LANDAU LIFSHITZ Quantum Mechanics Non-relativistic Theory Second edition, revised and enlarged Course of Theoretical Physics Volume 3 o £ 3 £ O 3 O V) Pergamon D. Landau and E. M. Lifshitz Institute of Physical Problems USSR Academy of Sciences Pergamon Press QUANTUM MECHANICS Non-relativistic Theory L D. LANDAU and E. M. UFSHITZ This second edition, published to meet the consistent demand for this important and informative book, has been considerably revised and enlarged, the basic plan and style of the first edition, however, being retained. The volume gives a comprehensive treatment of non-relativistic quantum mechanics, and an introduction to its application to atomic and molecular phenomena. Thetopics dealtwith include the basic concepts, Schrodinger's equation, angular momentum, motion in centrally symmetric fields perturbation theory, the quasi-classical case, spin, identity of particles, atoms, diatomic and polyatomic molecules, the theory of symmetry, elastic and inelastic collisions, and motion in a magnetic field. In this second edition, extensive changes have been made in the sections dealing with the theory of the addition of angular momenta and with collision theory and a new chapter on nuclear structure has been added. The discussion is intended to display the physical significance of the theory, and to be complete and self-contained. As with other volumes in this series, a list of which —together with statements aboutthem by knowledgeable and impartial reviewers — will be found on the back cover of this jacket, this book has been written to a level that will prove invaluable to those carrying out undergraduate and post-graduate study. 2217 COURSE OF THEORETICAL PHYSICS Volume 3 QUANTUM MECHANICS Non-relativistic Theory Second edition, revised and enlarged OTHER TITLES IN THE SERIES Vol.1. MECHANICS Vol. 2. THE CLASSICAL THEORY OF FIELDS Vol.3. QUANTUM MECHANICS— Non-Relativistic Theory Vol. 4. RELATIVISTIC PHYSICS Vol. 5. STATISTICAL PHYSICS Vol. 6. FLUID MECHANICS Vol. 7. THEORY OF ELASTICITY Vol. 8. ELECTRODYNAMICS OF CONTINUOUS MEDIA Vol. 9. PHYSICAL KINETICS QUANTUM MECHANICS NON-RELATIVISTIC THEORY by L. D. LANDAU and E. M. LIFSHITZ INSTITUTE OF PHYSICAL PROBLEMS, U.S.S.R. ACADEMY OF SCIENCES Volume 3 of Course of Theoretical Physics Translated from the Russian by J. B. SYKES and J. S. BELL Second edition, revised and enlarged PERGAMON PRESS OXFORD • LONDON • EDINBURGH • NEW YORK PARIS • FRANKFURT Pergamon Press Ltd., Headington Hill Hall, Oxford 4 & 5 Fitzroy Square, London W.l Pergamon Press (Scotland) Ltd., 2 & 3 Teviot Place, Edinburgh 1 Pergamon Press Inc., 44-01 21st Street, Long Island City, New York 11101 Pergamon Press S.A.R.L., 24 rue des Ecoles, Paris 5e Pergamon Press GmbH, Kaiserstrasse 75, Frankfurt-am-Main Sole distributors in the U.S.A. Addison- Wesley Publishing Company, Inc. Reading, Massachusetts Copyright © 1958 and 1965 Pergamon Press Ltd. First published in English 1958 2nd impression 1959 3rd impression 1962 Second {revised) edition 1965 Library of Congress Card Number 57-14444 Printed in Great Britain by J. W. Arrowsmith Ltd., Bristol. 2217/65 CONTENTS Page From the Preface to the first English edition xi Preface to the second English edition xii Notation xiii I. THE BASIC CONCEPTS OF QUANTUM MECHANICS §1. The uncertainty principle 1 §2. The principle of superposition 6 §3. Operators 8 §4. Addition and multiplication of operators 13 §5. The continuous spectrum 15 §6. The passage to the limiting case of classical mechanics 20 §7. The wave function and measurements 21 II. ENERGY AND MOMENTUM §8. The Hamiltonian operator 25 §9. The differentiation of operators with respect to time 26 §10. Stationary states 27 §11. Matrices 30 §12. Transformation of matrices 35 §13. The Heisenberg representation of operators 37 §14. The density matrix 38 §15. Momentum 41 §16. Uncertainty relations 46 III. SCHRODINGER'S EQUATION §17. Schrodinger's equation 50 §18. The fundamental properties of Schrodinger's equation 53 §19. The current density 55 §20. The variational principle 58 §21. General properties of motion in one dimension 60 §22. The potential well 63 §23. The linear oscillator 67 §24. Motion in a homogeneous field 73 §25. The transmission coefficient 75 IV. ANGULAR MOMENTUM §26. Angular momentum 81 §27. Eigenvalues of the angular momentum 85 §28. Eigenfunctions of the angular momentum 88 §29. Matrix elements of vectors 91 §30. Parity of a state 95 §31. Addition of angular momenta 97 vi Contents V. MOTION IN A CENTRALLY SYMMETRIC FIELD Page §32. Motion in a centrally symmetric field 101 §33. Free motion (spherical polar co-ordinates) 104 §34. Resolution of a plane wave 111 §35. "Fall" of a particle to the centre 113 §36. Motion in a Coulomb field (spherical polar co-ordinates) 116 §37. Motion in a Coulomb field (parabolic co-ordinates) 125 VI. PERTURBATION THEORY §38. Perturbations independent of time 129 §39. The secular equation 133 §40. Perturbations depending on time 136 §41. Transitions under a perturbation acting for a finite time 140 §42. Transitions under the action of a periodic perturbation 146 §43. Transitions in the continuous spectrum 147 §44. The uncertainty relation for energy 150 §45. Potential energy as a perturbation 153 VII. THE QUASI-CLASSICAL CASE §46. The wave function in the quasi-classical case 158 §47. Boundary conditions in the quasi- classical case 161 §48. Bohr and Sommerfeld's quantisation rule 162 §49. Quasi-classical motion in a centrally symmetric field 167 §50. Penetration through a potential barrier 171 §51. Calculation of the quasi-classical matrix elements 177 §52. The transition probability in the quasi-classical case 181 §53. Transitions under the action of adiabatic perturbations 185 VIII. SPIN §54. Spin 188 §55. Spinors 191 §56. Spinors of higher rank 196 §57. The wave functions of particles with arbitrary spin 198 §58. The relation between spinors and tensors 200 §59. Partial polarisation of particles 204 §60. Time reversal and Kramers' theorem 206 IX. IDENTITY OF PARTICLES §61. The principle of indistinguishability of similar particles 209 §62. Exchange interaction 212 §63. Symmetry with respect to interchange 216 §64. Second quantisation. The case of Bose statistics 221 §65. Second quantisation. The case of Fermi statistics 227 Contents vii X. THE ATOM Page §66. Atomic energy levels 231 §67. Electron states in the atom 232 §68. Hydrogen-like energy levels 236 §69. The self-consistent field 237 §70. The Thomas-Fermi equation 241 §71. Wave functions of the outer electrons near the nucleus 246 §72. Fine structure of atomic levels 247 §73. The periodic system of D. I. Mendeleev 252 §74. X-ray terms 259 §75. Multipole moments 261 §76. The Stark effect 265 §77. The Stark effect in hydrogen 269 XI. THE DIATOMIC MOLECULE §78. Electron terms in the diatomic molecule 277 §79. The intersection of electron terms 279 §80. The relation between molecular and atomic terms 282 §81. Valency 286 §82. Vibrational and rotational structures of singlet terms in the diatomic molecule 293 §83. Multiplet terms. Case a 299 §84. Multiplet terms. Case b 303 §85. Multiplet terms. Cases c and d 307 §86. Symmetry of molecular terms 309 §87. Matrix elements for the diatomic molecule 312 §88. A-doubling 316 §89. The interaction of atoms at large distances 319 §90. Pre-dissociation 322 XII. THE THEORY OF SYMMETRY §91. Symmetry transformations 332 §92. Transformation groups 335 §93. Point groups 338 §94. Representations of groups 347 §95. Irreducible representations of point groups 354 §96. Irreducible representations and the classification of terms 358 §97. Selection rules for matrix elements 361 §98. Continuous groups 364 §99. Two-valued representations of finite point groups 367 viii Contents XIII. POLYATOMIC MOLECULES Page §100. The classification of molecular vibrations 371 §101. Vibrational energy levels 378 §102. Stability of symmetrical configurations of the molecule 380 §103. Quantisation of the rotation of a rigid body 383 §104. The interaction between the vibrations and the rotation of the molecule 389 §105. The classification of molecular terms 394 XIV. ADDITION OF ANGULAR MOMENTA §106. 3/-symbols 401 §107. Matrix elements of tensors 408 §108. 6/-symbols 412 §109. Matrix elements for addition of angular momenta 418 XV. MOTION IN A MAGNETIC FIELD §110. Schrodinger's equation in a magnetic field 421 §111. Motion in a uniform magnetic field 424 §112. The Zeeman effect 427 §113. Spin in a variable magnetic field 434 §114. The current density in a magnetic field 435 XVI. NUCLEAR STRUCTURE §115. Isotopic invariance 438 §116. Nuclear forces 442 §117. The shell model 447 §118. Non-spherical nuclei 456 §119. Isotopic shift 461 §120. Hyperfine structure of atomic levels 463 §121. Hyperfine structure of molecular levels 466 XVII. THE THEORY OF ELASTIC COLLISIONS §122. The general theory of scattering 469 §123. An investigation of the general formula 472 §124. The unitary condition for scattering 475 §125. Born's formula 479 §126. The quasi-classical case 486 §127. Scattering at high energies 489 §128. Analytical properties of the scattering amplitude 492 §129. The dispersion relation 497 §130. The scattering of slow particles 500 Contents ix Page §131. Resonance scattering at low energies 505 §132. Resonance at a quasi-discrete level 511 §133. Rutherford's formula 516 §134. The system of wave functions of the continuous spectrum 519 §135. Collisions of like particles 523 §136. Resonance scattering of charged particles 526 §137. Elastic collisions between fast electrons and atoms 531 §138. Scattering with spin-orbit interaction 535 XVIII. THE THEORY OF INELASTIC COLLISIONS §139. Elastic scattering in the presence of inelastic processes 542 §140. Inelastic scattering of slow particles 548 §141. The scattering matrix in the presence of reactions 550 §142. Breit and Wigner's formula 554 §143. Interaction in the final state in reactions 562 §144. Behaviour of cross-sections near the reaction threshold 565 §145. Inelastic collisions between fast electrons and atoms 571 §146. The effective retardation 580 §147. Inelastic collisions between heavy particles and atoms 584 §148. Scattering by molecules 587 MATHEMATICAL APPENDICES §a. Hermite polynomials 593 §b. The Airy function 596 §c. Legendre polynomials 598 §d. The confluent hypergeometric function 600 §e. The hypergeometric function 605 §f. The calculation of integrals containing confluent hypergeometric functions ■ 607 INDEX 611 FROM THE PREFACE TO THE FIRST ENGLISH EDITION The present book is one of the series on Theoretical Physics, in which we endeavour to give an up-to-date account of various departments of that science. The complete series will contain the following nine volumes : 1. Mechanics. 2. The classical theory of fields. 3. Quantum mechanics (non-relativistic theory). 4. Relativistic quantum theory. 5. Statistical physics. 6. Fluid mechanics. 7. Theory of elasticity. 8. Electrodynamics of continuous media. 9. Physical kinetics. Of these, volumes 4 and 9 remain to be written. The scope of modern theoretical physics is very wide, and we have, of course, made no attempt to discuss in these books all that is now included in the subject. One of the principles which guided our choice of material was not to deal with those topics which could not properly be expounded without at the same time giving a detailed account of the existing experimental results. For this reason the greater part of nuclear physics, for example, lies outside the scope of these books. Another principle of selection was not to discuss very complicated applications of the theory. Both these criteria are, of course, to some extent subjective. We have tried to deal as fully as possible with those topics that are included. For this reason we do not, as a rule, give references to the original papers, but simply name their authors. We give bibliographical references only to work which contains matters not fully expounded by us, which by their com- plexity lie "on the borderline" as regards selection or rejection. We have tried also to indicate sources of material which might be of use for reference. Even with these limitations, however, the bibliography given makes no pre- tence of being exhaustive. We attempt to discuss general topics in such a way that the physical signifi- cance of the theory is exhibited as clearly as possible, and then to build up the mathematical formalism. In doing so, we do not aim at "mathematical rigour" of exposition, which in theoretical physics often amounts to self- deception. The present volume is devoted to non-relativistic quantum mechanics. By "relativistic theory" we here mean, in the widest sense, the theory of all quantum phenomena which significantly depend on the velocity of light. The volume on this subject (volume 4) will therefore contain not only Dirac's relativistic theory and what is now known as quantum electrodynamics, but also the whole of the quantum theory of radiation. Institute of Physical Problems L. D. Landau USSR Academy of Sciences E. M. Lifshitz August 1956 PREFACE TO THE SECOND ENGLISH EDITION For this second edition the book has been considerably revised and en- larged, but the general plan and style remain as before. Every chapter has been revised. In particular, extensive changes have been made in the sections dealing with the theory of the addition of angular momenta and with collision theory. A new chapter on nuclear structure has been added; in accordance with the general plan of the course, the subjects in question are discussed only to the extent that is proper without an accompanying detailed analysis of the experimental results. We should like to express our thanks to all our many colleagues whose comments have been utilised in the revision of the book. Numerous com- ments were received from V. L. Ginzburg and Ya. A. Smorodinskii. We are especially grateful to L. P. Pitaevskii for the great help which he has given in checking the formulae and the problems. Our sincere thanks are due to Dr. Sykes and Dr. Bell, who not only translated excellently both the first and the second edition of the book, but also made a number of useful comments and assisted in the detection of various misprints in the first edition. Finally, we are grateful to the Pergamon Press, which always acceded to our requests during the production of the book. L. D. Landau October 1964 E. M. Lifshitz NOTATION Operators are denoted by a circumflex dq element in configuration space f nm = /" = (n\f\m) matrix elements of the quantity/ (see definition in §11) % m = (E n — Em)!^ transition frequency {/, g} = fg — g/ commutator of two operators i? Hamiltonian S, &? electric and magnetic fields Si phase shifts of wave functions eua antisymmetric unit tensor ■"± = "a; i ^"y CHAPTER I THE BASIC CONCEPTS OF QUANTUM MECHANICS §1. The uncertainty principle When we attempt to apply classical mechanics and electrodynamics to explain atomic phenomena, they lead to results which are in obvious conflict with experiment. This is very clearly seen from the contradiction obtained on applying ordinary electrodynamics to a model of an atom in which the elec- trons move round the nucleus in classical orbits. During such motion, as in any accelerated motion of charges, the electrons would have to emit electro- magnetic waves continually. By this emission, the electrons would lose their energy, and this would eventually cause them to fall into the nucleus. Thus, according to classical electrodynamics, the atom would be unstable, which does not at all agree with reality. This marked contradiction between theory and experiment indicates that the construction of a theory applicable to atomic phenomena — that is, pheno- mena occurring in particles of very small mass at very small distances — demands a fundamental modification of the basic physical concepts and laws. As a starting-point for an investigation of these modifications, it is conveni- ent to take the experimentally observed phenomenon known as electron diffraction.^ It is found that, when a homogeneous beam of electrons passes through a crystal, the emergent beam exhibits a pattern of alternate maxima and minima of intensity, wholly similar to the diffraction pattern observed in the diffraction of electromagnetic waves. Thus, under certain conditions, the behaviour of material particles — in this case, the electrons — displays features belonging to wave processes. How markedly this phenomenon contradicts the usual ideas of motion is best seen from the following imaginary experiment, an idealisation of the experiment of electron diffraction by a crystal. Let us imagine a screen impermeable to electrons, in which two slits are cut. On observing the passage of a beam of electrons^ through one of the slits, the other being covered, we obtain, on a continuous screen placed behind the slit, some pat- tern of intensity distribution; in the same way, by uncovering the second slit and covering the first, we obtain another pattern. On observing the passage of the beam through both slits, we should expect, on the basis of ordinary classical ideas, a pattern which is a simple superposition of the other two : each electron, moving in its path, passes through one of the slits and f The phenomenon of electron diffraction was in fact discovered after quantum mechanics was invented. In our discussion, however, we shall not adhere to the historical sequence of development of the theory, but shall endeavour to construct it in such a way that the connection between the basic principles of quantum mechanics and the experimentally observed phenomena is most clearly shown J The beam is supposed so rarefied that the interaction of the particles in it plays no part. 1 2 The Basic Concepts of Quantum Mechanics §1 has no effect on the electrons passing through the other slit. The phenomenon of electron diffraction shows, however, that in reality we obtain a diffraction pattern which, owing to interference, does not at all correspond to the sum of the patterns given by each slit separately. It is clear that this result can in no way be reconciled with the idea that electrons move in paths. Thus the mechanics which governs atomic phenomena — quantum mechanics or wave mechanics — must be based on ideas of motion which are fundamentally different from those of classical mechanics. In quantum mechanics there is no such concept as the path of a particle. This forms the content of what is called the uncertainty principle, one of the fundamental principles of quantum mechanics, discovered by W. Heisenberg in 1927.f In that it rejects the ordinary ideas of classical mechanics, the uncertainty principle might be said to be negative in content. Of course, this principle in itself does not suffice as a basis on which to construct a new mechanics of particles. Such a theory must naturally be founded on some positive asser- tions, which we shall discuss below (§2). However, in order to formulate these assertions, we must first ascertain the statement of the problems which confront quantum mechanics. To do so, we first examine the special nature of the interrelation between quantum mechanics and classical mechanics. A more general theory can usually be formulated in a logically complete manner, independently of a less general theory which forms a limiting case of it. Thus, relativistic mechanics can be constructed on the basis of its own fundamental principles, without any reference to Newtonian mechanics. It is in principle impossible, however, to formulate the basic concepts of quantum mechanics without using classical mechanics. The fact that an electron^ has no definite path means that it has also, in itself, no other dynamical characteristics. || Hence it is clear that, for a system composed only of quantum objects, it would be entirely impossible to construct any logically independent mechanics. The possibility of a quantitative description of the motion of an electron requires the presence also of physical objects which obey classical mechanics to a sufficient degree of accuracy. If an electron interacts with such a "classical object", the state of the latter is, generally speaking, altered. The nature and magnitude of this change depend on the state of the electron, and therefore may serve to characterise it quantitatively. In this connection the "classical object" is usually called apparatus, and its interaction with the electron is spoken of as measurement. However, it must be emphasised that we are here not discussing a process of measurement in which the physicist-observer takes part. By measurement, in quantum mechanics, we understand any process of interaction between classical and f It is of interest to note that the complete mathematical formalism of quantum mechanics was constructed by W. Heisenberg and E. Schrodinger in 1925-6, before the discovery of the uncertainty principle, which revealed the physical content of this formalism. X In this and the following sections we shall, for brevity, speak of "an electron", meaning in general any object of a quantum nature, i.e. a particle or system of particles obeying quantum mechanics and not classical mechanics. I We refer to quantities which characterise the motion of the electron, and not to those, such as the charge and the mass, which relate to it as a particle ; these are parameters. §1 The uncertainty principle 3 quantum objects, occurring apart from and independently of any observer. The importance of the concept of measurement in quantum mechanics was elucidated by N. Bohr. We have denned "apparatus" as a physical object which is governed, with sufficient accuracy, by classical mechanics. Such, for instance, is a body of large enough mass. However, it must not be supposed that apparatus is necessarily macroscopic. Under certain conditions, the part of apparatus may also be taken by an object which is microscopic, since the idea of "with sufficient accuracy" depends on the actual problem proposed. Thus, the motion of an electron in a Wilson chamber is observed by means of the cloudy track which it leaves, and the thickness of this is large compared with atomic dimensions; when the path is determined with such low accuracy, the electron is an entirely classical object. Thus quantum mechanics occupies a very unusual place among physical theories: it contains classical mechanics as a limiting case, yet at the same time it requires this limiting case for its own formulation. We may now formulate the problem of quantum mechanics. A typical problem consists in predicting the result of a subsequent measurement from the known results of previous measurements. Moreover, we shall see later that, in comparison with classical mechanics, quantum mechanics, generally speaking, restricts the range of values which can be taken by various physical quantities (for example, energy) : that is, the values which can be obtained as a result of measuring the quantity concerned. The methods of quantum mechanics must enable us to determine these admissible values. The measuring process has in quantum mechanics a very important pro- perty: it always affects the electron subjected to it, and it is in principle impossible to make its effect arbitrarily small, for a given accuracy of measure- ment. The more exact the measurement, the stronger the effect exerted by it, and only in measurements of very low accuracy can the effect on the mea- sured object be small. This property of measurements is logically related to the fact that the dynamical characteristics of the electron appear only as a result of the measurement itself. It is clear that, if the effect of the measuring process on the object of it could be made arbitrarily small, this would mean that the measured quantity has in itself a definite value independent of the measurement. Among the various kinds of measurement, the measurement of the co- ordinates of the electron plays a fundamental part. Within the limits of applicability of quantum mechanics, a measurement of the co-ordinates of an electron can always be performed-]- with any desired accuracy. Let us suppose that, at definite time intervals At, successive measurements of the co-ordinates of an electron are made. The results will not in general lie on a smooth curve. On the contrary, the more accurately the measurements f Once again we emphasise that, in speaking of "performing a measurement", we refer to the interaction of an electron with a classical " apparatus", which in no way presupposes the presence of an external observer. 4 The Basic Concepts of Quantum Mechanics §1 are made, the more discontinuous and disorderly will be the variation of their results, in accordance with the non-existence of a path of the electron. A fairly smooth path is obtained only if the co-ordinates of the electron are measured with a low degree of accuracy, as for instance from the condensa- tion of vapour droplets in a Wilson chamber. If now, leaving the accuracy of the measurements unchanged, we diminish the intervals At between measurements, then adjacent measurements, of course, give neighbouring values of the co-ordinates. However, the results of a series of successive measurements, though they lie in a small region of space, will be distributed in this region in a wholly irregular manner, lying on no smooth curve. In particular, as A* tends to zero, the results of adjacent measurements by no means tend to lie on one straight line. This circumstance shows that, in quantum mechanics, there is no such concept as the velocity of a particle in the classical sense of the word, i.e. the limit to which the difference of the co-ordinates at two instants, divided by the interval At between these instants, tends as At tends to zero. However, we shall see laier that in quantum mechanics, nevertheless, a reasonable definition of the velocity of a particle at a given instant can be constructed, and this velocity passes into the classical velocity as we pass to classical mech- anics. But whereas in classical mechanics a particle has definite co-ordinates and velocity at any given instant, in quantum mechanics the situation is entirely different. If, as a result of measurement, the electron is found to have definite co-ordinates, then it has no definite velocity whatever. Conversely, if the electron has a definite velocity, it cannot have a definite position in space. For the simultaneous existence of the co-ordinates and velocity would mean the existence of a definite path, which the electron has not. Thus, in quantum mechanics, the co-ordinates and velocity of an electron are quantities which cannot be simultaneously measured exactly, i.e. they cannot simultane- ously have definite values. We may say that the co-ordinates and velocity of the electron are quantities which do not exist simultaneously. In what follows we shall derive the quantitative relation which determines the pos- sibility of an inexact measurement of the co-ordinates and velocity at the same instant. A complete description of the state of a physical system in classical mech- anics is effected by stating all its co-ordinates and velocities at a given instant ; with these initial data, the equations of motion completely determine the behaviour of the system at all subsequent instants. In quantum mechanics such a description is in principle impossible, since the co-ordinates and the corresponding velocities cannot exist simultaneously. Thus a description of the state of a quantum system is effected by means of a smaller number of quantities than in classical mechanics, i.e. it is less detailed than a classical description. A very important consequence follows from this regarding the nature of the predictions made in quantum mechanics. Whereas a classical description suffices to predict the future motion of a mechanical system with complete §1 The uncertainty principle 5 accuracy, the less detailed description given in quantum mechanics evidently cannot be enough to do this. This means that, even if an electron is in a state described in the most complete manner possible in quantum mechanics, its behaviour at subsequent instants is still in principle uncertain. Hence quan- tum mechanics cannot make completely definite predictions concerning the future behaviour of the electron. For a given initial state of the electron, a subsequent measurement can give various results. The problem in quantum mechanics consists in determining the probability of obtaining vari- ous results on performing this measurement. It is understood, of course, that in some cases the probability of a given result of measurement may be equal to unity, i.e. certainty, so that the result of that measurement is unique. All measuring processes in quantum mechanics may be divided into two classes. In one, which contains the majority of measurements, we find those which do not, in any state of the system, lead with certainty to a unique result. The other class contains measurements such that for every possible result of measurement there is a state in which the measurement leads with certainty to that result. These latter measurements, which may be called predictable, play an important part in quantum mechanics. The quantitative characteristics of a state which are determined by such measurements are what are called physical quantities in quantum mechanics. If in some state a measurement gives with certainty a unique result, we shall say that in this state the corresponding physical quantity has a definite value. In future we shall always understand the expression "physical quantity" in the sense given here. We shall often find in what follows that by no means every set of physical quantities in quantum mechanics can be measured simultaneously, i.e. can all have definite values at the same time. We have already mentioned one example, namely the velocity and co-ordinates of an electron. An important part is played in quantum mechanics by sets of physical quantities having the following property: these quantities can be measured simultaneously, but if they simultaneously have definite values, no other physical quantity (not being a function of these) can have a definite value in that state. We shall speak of such sets of physical quantities as complete sets; in particular cases a complete set may consist of only one quantity. Any description of the state of an electron arises as a result of some mea- surement. We shall now formulate the meaning of a complete description of a state in quantum mechanics. Completely described states occur as a result of the simultaneous measurement of a complete set of physical quanti- ties. From the results of such a measurement we can, in particular, deter- mine the probability of various results of any subsequent measurement, regardless of the history of the electron prior to the first measurement. In quantum mechanics we need concern ourselves in practice only with completely described states, and from now on (except in §14) we shall under- stand by the states of a quantum system just these completely described states. 6 The Basic Concepts of Quantum Mechanics §2 §2. The principle of superposition Passing now to an exposition of the fundamental mathematical formalism of quantum mechanics, we shall denote by q the set of co-ordinates of a quan- tum system, and by dq the product of the differentials of these co-ordinates. This dq is often called an element of volume in the configuration space of the system; for one particle, dq coincides with an element of volume dV in ordinary space. The basis of the mathematical formalism of quantum mechanics lies in the fact that any state of a system can be described, at a given moment, by a definite (in general complex) function Y(#) of the co-ordinates. The square of the modulus of this function determines the probability distribution of the values of the co-ordinates: |Y| 2 d# is the probability that a measurement performed on the system will find the values of the co-ordinates to be in the element dq of configuration space. The function Y is called the wave function of the system (sometimes also the probability amplitude), j- A knowledge of the wave function allows us, in principle, to calculate the probability of the various results of any other measurement (not of the co- ordinates) also. All these probabilities are determined by expressions bi- linear in Y and Y*. The most general form of such an expression is jjV{q)Wffl<Kq t <t)*qM, (2.1) where the function (f>(q, q) depends on the nature and the result of the mea- surement, and the integration is extended over all configuration space. The probability YY* of various values of the co-ordinates is itself an expression of this type. J The state of the system, and with it the wave function, in general varies with time. In this sense the wave function can be regarded as a function of time also. If the wave function is known at some initial instant, then, from the very meaning of the concept of complete description of a state, it is in principle determined at every succeeding instant. The actual dependence of the wave function on time is determined by equations which will be de- rived later. The sum of the probabilities of all possible values of the co-ordinates of the system must, by definition, be equal to unity. It is therefore necessary that the result of integrating |Y| 2 over all configuration space should be equal to unity: J>|»df = l. (2.2) This equation is what is called the normalisation condition for wave functions. If the integral of |Y| 2 converges, then by choosing an appropriate constant coefficient the function Y can always be, as we say, normalised. Sometimes, f It was first introduced into quantum mechanics by Schrodinger in 1926. % It is obtained from (2.1) when<£(g, q') = 8(?-? ) 8(q'-q ), where 8 denotes the delta function, defined in §5 below; q denotes the value of the co-ordinates whose probability is required. §2 The principle of superposition 7 however, wave functions are used which are not normalised; moreover, we shall see later that the integral of |TJ 2 may diverge, and then Y cannot be normalised by the condition (2.2). In such cases |T| 2 does not, of course, determine the absolute values of the probability of the co-ordinates, but the ratio of the values of |T| 2 at two different points of configuration space deter- mines the relative probability of the corresponding values of the co-ordinates. Since all quantities calculated by means of the wave function, and having a direct physical meaning, are of the form (2.1), in which T appears multiplied by Y*, it is clear that the normalised wave function is determined only to within a constant phase fact or of the form e ia (where a is any real number), whose modulus is unity. This indeterminacy is in principle irremovable; it is, however, unimportant, since it has no effect upon any physical results. The positive content of quantum mechanics is founded on a series of propositions concerning the properties of the wave function. These are as follows. Suppose that, in a state with wave function VF^), some measurement leads with certainty to a definite result (result 1), while in a state with Y 2 (#) it leads to result 2. Then it is assumed that every linear combination of Y x and Y 2 , i.e. every function of the form c J T 1 +c 2 x F 2 (where c x and c % me con- stants), gives a state in which that measurement leads to either result 1 or result 2. Moreover, we can assert that, if we know the time dependence of the states, which for the one case is given by the function T^, t), and for the other by Y 2 (<7, t), then any linear combination also gives a possible dependence of a state on time. These propositions can be immediately generalised to any number of different states. The above set of assertions regarding wave functions constitutes what is called the principle of superposition of states, the chief positive principle of quantum mechanics. In particular, it follows at once from this principle that all equations satisfied by wave functions must be linear in Y. Let us consider a system composed of two parts, and suppose that the state of this system is given in such a way that each of its parts is completely described.f Then we can say that the probabilities of the co-ordinates q x of the first part are independent of the probabilities of the co-ordinates q % of the second part, and therefore the probability distribution for the whole system should be equal to the product of the probabilities of its parts. This means that the wave function Y 12 (ft, q z ) of the system can be represented in the form of a product of the wave functions Y^ft) and Y 2 (# 2 ) of its parts: ^i 2 (?i,? 2 )=Yi(?i)Y 2 (? 2 ). (2.3) If the two parts do not interact, then this relation between the wave function of the system and those of its parts will be maintained at future instants also, f This, of course, means that the state of the whole system is completely described also. However, we emphasise that the converse statement is by no means true: a complete description of the state of the whole system does not in general completely determine the states of its individual parts (see also §14). 8 The Basic Concepts of Quantum Mechanics §3 i.e. we can write Tu(fo ?i, t) = YjCfc, /) T 2 ( ?2 , 0- (2.4) §3. Operators Let us consider some physical quantity / which characterises the state of a quantum system. Strictly, we should speak in the following discussion not of one quantity, but of a complete set of them at the same time. However, the discussion is not essentially changed by this, and for brevity and simplicity we shall work below in terms of only one physical quantity. The values which a given physical quantity can take are called in quantum mechanics its eigenvalues, and the set of these is referred to as the spectrum of eigenvalues of the given quantity. In classical mechanics, generally speak- ing, quantities run through a continuous series of values. In quantum mech- anics also there are physical quantities (for instance, the co-ordinates) whose eigenvalues occupy a continuous range ; in such cases we speak of a continuous spectrum of eigenvalues. As well as such quantities, however, there exist in quantum mechanics others whose eigenvalues form some discrete set; in such cases we speak of a discrete spectrum. We shall suppose for simplicity that the quantity / considered here has a discrete spectrum; the case of a continuous spectrum will be discussed in §5. The eigenvalues of the quantity / are denoted by/ n , where the suffix n takes the values 0, 1, 2, 3 We also denote the wave function of the system, in the state where the quantity / has the value / n , by Y n . The wave functions Y n are called the eigenfunctions of the given physical quantity/. Each of these functions is supposed normalised, so that J> n |*d«z = l. (3.1) If the system is in some arbitrary state with wave function T, a measure- ment of the quantity / carried out on it will give as a result one of the eigen- values f n . In accordance with the principle of superposition, we can assert that the wave function Y must be a linear combination of those eigenfunc- tions T TC which correspond to the values f n that can be obtained, with prob- ability different from zero, when a measurement is made on the system and it is in the state considered. Hence, in the general case of an arbitrary state, the function T can be represented in the form of a series Y=Sa w Y w , (3.2) where the summation extends over all n, and the a n are some constant coeffi- cients. Thus we reach the conclusion that any wave function can be, as we say, expanded in terms of the eigenfunctions of any physical quantity. A set of functions in terms of which such an expansion can be made is called a complete (or closed) set. §3 Operators 9 The expansion (3.2) makes it possible to determine the probability of find- ing (i.e. the probability of getting the corresponding result on measurement), in a system in a state with wave function Y, any given value/ n of the quantity /. For, according to what was said in the previous section, these probabili- ties must be determined by some expressions bilinear in Y and Y*, and therefore must be bilinear in a n and a n *. Furthermore, these expressions must, of course, be positive. Finally, the probability of the value f n must become unity if the system is in a state with wave function Y = Y n , and must become zero if there is no term containing Y n in the expansion (3.2) of the wave function Y. This means that the required probability must be unity if all the coefficients a n except one (with the given n) are zero, that one being unity; the probability must be zero, if the a n concerned is zero. The only essentially positive quantity satisfying these conditions is the square of the modulus of the coefficient a n . Thus we reach the result that the squared modulus \a n \ % of each coefficient in the expansion (3.2) determines the prob- ability of the corresponding value f n of the quantity / in the state with wave function Y. The sum of the probabilities of all possible values f n must be equal to unity; in other words, the relation S la w | 2 = 1 (3.3) n must hold. If the function Y were not normalised, then the relation (3.3) would not hold either. The sum £ \a n \ 2 would then be given by some expression bilinear in Y and Y*, and becoming unity when Y was normalised. Only the integral J YY* dq is such an expression. Thus the equation SflA* = JYY*d ? (3.4) must hold. On the other hand, multiplying by Y the expansion Y* = S tf n *Y n * of the function Y* (the complex conjugate of Y), and integrating, we obtain j YY* dq = S a n *j Y n *Y dq. Comparing this with (3.4), we have Sa B « tt * = Sa n *jY n *Yd ? , from which we derive the following formula determining the coefficients a n in the expansion of the function Y in terms of the eigenfunctions Y n : a n = jYY n *d 2 . (3.5) If we substitute here from (3.2), we obtain « n = Sa w fY TO Y n *d 9 , 10 The Basic Concepts of Quantum Mechanics §3 from which it is evident that the eigenfunctions must satisfy the conditions JT m V n *dq = 8 nm , (3.6) where $ nm = 1 for n = m and S nm = for n # m. The fact that the integrals of the products T w T n * with m ^ n vanish is called the orthogonality of the functions T n . Thus the set of eigenfunctions T n forms a complete set of normalised and orthogonal (or, for brevity, orthonormal) functions. We shall now introduce the concept of the mean value f of the quantity/ in the given state. In accordance with the usual definition of mean values, we define / as the sum of all the eigenvalues f n of the given quantity, each multiplied by the corresponding probability |a n | 2 . Thus /= IfnW- (3.7) We shall write / in the form of an expression which does not contain the coefficients a n in the expansion of the function T, but this function itself. Since the products a n a n * appear in (3.7), it is clear that the required expres- sion must be bilinear in T and Y*. We introduce a mathematical opera- tor, which we denotef by /and define as follows. Let (/Y) denote the result of the operator / acting on the function Y. We define / in such a way that the integral of the product of (/Y) and the complex conjugate function Y* is equal to the mean value /: /=jV(/T)d ? . (3.8) It is easily seen that, in the general case, the operator / is a linear J integral operator. For, using the expression (3.5) for a n , we can rewrite the definition (3.7) of the mean value in the form /= \f n a n a* = J Y*(S a n f n W n ) 6q. Comparing this with (3.8), we see that the result of the operator / acting on the function Y has the form (/Y) = Sa n / M Y w . (3.9) If we substitute here the expression (3.5) for a n , we find that /is an integral operator of the form tfr) = JK(q,<nv(<nw, (3.io) where the function K(q, q') (called the kernel of the operator) is K(q, q') = S/.Y.-foT^fe). (3.11) f By convention, we shall always denote operators by letters with circumflexes. + An operator is said to be linear if it has the properties /(^i+T 2 ) =M+/Y, and/(«T) = afY, where X F 1 and *F 2 are arbitrary functions and a is an arbitrary constant. §3 Operators 11 Thus, for every physical quantity in quantum mechanics, there is a definite corresponding linear operator. It is seen from (3.9) that, if the function T is one of the eigenfunctions \F n (so that all the a n except one are zero), then, when the operator / acts on it, this function is simply multiplied by the corresponding eigenvalue f n : fa =/«^n- (3.12) (In what follows we shall always omit the parentheses in the expression (/T), where this cannot cause any misunderstanding; the operator is taken to act on the expression which follows it.) Thus we can say that the eigen- functions of the given physical quantity /are the solutions of the equation where / is a constant, and the eigenvalues are the values of this constant for which the above equation has solutions satisfying the required conditions. Of course, while the operator /is still defined only by the expressions (3.10) and (3.11), which themselves contain the eigenfunctions ^I^, no further con- clusions can be drawn from the result we have obtained. However, as we shall see below, the form of the operators for various physical quantities can be determined from direct physical considerations, and then the above pro- perty of the operators enables us to find the eigenfunctions and eigenvalues by solving the equations p¥ = f¥. The values which can be taken by real physical quantities are obviously real. Hence the mean value of a physical quantity must also be real, in any state. Conversely, if the mean value of a physical quantity is real in every state, its eigenvalues also are all real; to show this, it is sufficient to note that the mean values coincide with the eigenvalues in the states described by the functions T n . From the fact that the mean values are real, we can draw some conclusions concerning the properties of operators. Equating the expression (3.8) to its complex conjugate, we obtain the relation J T*(/Y) dg = j Y(f*Y*) dg, (3.13) where /* denotes the operator which is the complex conjugate of /. j- This relation does not hold in general for an arbitrary linear operator, so that it is a restriction on the form of the operator /. For an arbitrary operator / we can find what is called the transposed operator/, defined in such a way that JY(fl>)dq=f<l>(jY)dq, (3.14) where *F and <D are two different functions. If we take, as the function O, the function T* which is the complex conjugate of Y, then a comparison with (3.13) shows that we must have /=/•• (3.15) t By definition, if for^the operator /we h&vefifi = <f>, then the complex conjugate operator /* is that for which we have /*^t* = <f>*. 12 The Basic Concepts of Quantum Mechanics §3 Operators satisfying this condition are said to be Hermitian.^ Thus the operators corresponding, in the mathematical formalism of quantum mechanics, to real physical quantities must be Hermitian. We can formally consider complex physical quantities also, i.e. those whose eigenvalues are complex. Let / be such a quantity. Then we can introduce its complex conjugate quantity/*, whose eigenvalues are the com- plex conjugates of those of/. We denote by /+ the operator corresponding to the quantity/*. It is called the Hermitian conjugate of the operator / and, in general, will be different from the complex conjugate operator /* : from the condition/* = (/)* we find at once that / + =7*. (3.16) from which it is clear that /+ is in general not the same as /*. For a real physical quantity / = /+, i.e. the operator is the same as its Hermitian conjugate (Hermitian operators are also called self -conjugate). We shall show how the orthogonality of the eigenfunctions of an Hermitian operator corresponding to different eigenvalues can be directly proved. Let f n and/„, be two different eigenvalues of the quantity/, and T n , *F m the cor- responding eigenfunctions : f^n^fn^n, f X ¥ m =fm X r m - Multiplying both sides of the first of these equations by T m *, and both sides of the complex conjugate of the second by T n , and subtracting corre- sponding terms, we find W *fw _tF /*tF * — ( f —f W* W * We integrate both sides of this equation over q. Since /* =/, by (3.14) the integral on the left-hand side of the equation is zero, so that we have (fn-f m ) j T n Y w *d^ = 0, whence, since f n # f m , we obtain the required orthogonality property of the functions x P n and T m . We have spoken here of only one physical quantity /, whereas, as we said at the beginning of this section, we should have spoken of a complete set of physical quantities. We should then have found that to each of these quantities/ g, ... there corresponds its operator /, g, ... . The eigenfunctions T n then correspond to states in which all the quantities concerned have definite values, i.e. they correspond to definite sets of eigenvalues f n , g n , ... , and are simultaneous solutions of the system of equations /F=/V, iW=gY,.... f For a linear integral operator of the form (3.10), the Hermitian condition means that the kernel of the operator must be such that K(q, q') = K*(q', q). §4 Addition and multiplication of operators 13 §4. Addition and multiplication of operators Let / and g be two physical quantities which can simultaneously take definite values, and / and g their operators. The eigenvalues of the sum f+g of these quantities are equal to the sums of the eigenvalues of/ and g. To this new quantity f+g there will obviously correspond an operator equal to the sum of the operators / and g. For, if Y n are the eigenfunctions com- mon to the operators / and g } then it follows from /*F n = f^¥ ni £*F n =g n x ^n that (f+gWn = (fn+gnWn, i.e. the eigenvalues of the operator f+j> are equal to the sums/ n +£ n « If the quantities / and g cannot simultaneously take definite values, then it is meaningless to speak of their sum in the direct sense just mentioned. It is conventional in quantum mechanics to define the sum of the quantities / and g in such cases as the quantity whose mean value in an arbitrary state is equal to the sum of the mean values /and g: 7+J=/+|. (4-1) It is clear that, to the quantity f+g so defined, there corresponds an operator j+£. For, by formula (3.8), we have 7+g = \^*{f+gW d? = J>*/*F dq+ JVfF dq =f+g. The eigenvalues and eigenfunctions of the operator f+g will not, in general* now bear any relation to those of the quantities / and g. It is evident that, if the operators / and g are self-conjugate, the operator f+g will be so too, so that its eigenvalues are real and are those of the new quantity f+g thus defined. The following theorem should be noted. Let / and g be the smallest eigenvalues of the quantities / and g, and (f+g) that of the quantity f+g. Then (f+g)» >fo+go- (4.2) The equality holds if / and g can be measured simultaneously. The proof follows from the obvious fact that the mean value of a quantity is always greater than or equal to its least eigenvalue. In a state in which the quantity f+g has the value (f+g) we have f+g = (f+g) , and since, on the other hand, f+g = f+g ^ f +g > we arrive at the inequality (4.2). Next, let/ and g once more be quantities that can be measured simultane- ously. Besides their sum, we can also introduce the concept of their product as being a quantity whose eigenvalues are equal to the products of those of the quantities / and g. It is easy to see that, to this quantity, there corresponds an operator whose effect consists of the successive action on the function of first one and then the other operator. Such an operator is represented mathematically by the product of the operators / and §. For, if T n are the 14 The Basic Concepts of Quantum Mechanics §4 eigenfunctions common to the operators / and g, we have Jg?n =JtpTJ = fgn^n = gnf^n = gnfn^n (the symbol fg denotes an operator whose effect on a function ¥ consists of the successive action first of the operator g on the function ¥ and then of the operator / on the function #¥). We could equally well take the operator gf instead of fg, the former differing from the latter in the order of its factors. It is obvious that the result of the action of either of these operators on the functions ¥ n will be the same. Since, however, every wave function ¥ can be represented as a linear combination of the functions ¥ TC , it follows that the result of the action of the operators fg and gf on an arbitrary function will also be the same. This fact can be written in the form of the symbolic equation fg = gf or /g-gf=0. (4.3) Two such operators / and g are said to commute with each other. Thus we arrive at the important result : if two quantities/ and g can simultaneously take definite values, then their operators commute with each other. The converse theorem can also be proved (§11): if the operators /and g commute, then all their eigenfunctions can be taken common to both; physically, this means that the corresponding physical quantities can be measured simultaneously. Thus the commutability of the operators is a necessary and sufficient condition for the physical quantities to be simultane- ously measurable. A particular case of the product of operators is an operator raised to some power. From the above discussion we can deduce that the eigenvalues of an operator f p (where p is an integer) are equal to the pth. powers of the eigen- values of the operator/. Any function </>(/) of an operator can be defined as an operator whose eigenvalues are equal to the same function <£(/) of the eigenvalues of the operator/. If the function <f>(f) can be expanded as a Taylor series, this expresses the effect of the operator <£( / ) in terms of those of various powers/^. In particular, the operator / _1 is called the inverse of the operator / It is evident that the successive action of the operators / and / _1 on any function leaves the latter unchanged, i.e. // -1 = / _1 / = 1- If the quantities / and g cannot simultaneously take definite values, the concept of their product cannot be defined in the above manner. This appears in the fact that the operator fg is not self-conjugate in this case, and hence cannot correspond to any physical quantity. For, by the definition of the transpose of an operator we can write J" Yjg* dq = j¥/(i<D) dq = j (£<D)(/¥) dq. Here the operator/ acts only on the function ¥, and the operator £ on <E>, so that the integrand is a simple product of two functions^® and/¥. Again §5 The continuous spectrum 15 using the definition of the transpose of an operator, we can write Thus we obtain an integral in which the functions *F and <S> have changed places as compared with the original one. In other words, the operator §f is the transpose of fg, and we can write fg=if, (4-4) i.e. the transpose of the product fg is the product of the transposes of the factors written in the opposite order. Taking the complex conjugate of both sides of equation (4.4), we have (fg) + = g + f + - (4.5) If each of the operators / and g is Hermitian, then ( fg) + = gf It follows from this that the operator fg is Hermitian if and only if the factors / and g commute. We note that, from the products)^ and gf of two non-commuting Hermitian operators, we can form an Hermitian operator by taking the symmetrical combination iifg+gf)- (4.6) Such expressions are sometimes needed ; they are called symmetrised pro- ducts. It is easy to see that the difference fg—gf is an anti-Hermitian operator (i.e. one for which the transpose is equal to the complex conjugate taken with the opposite sign). It can be made Hermitian by multiplying by i; thus i(fg-gf) is again an Hermitian operator. In what follows we shall sometimes use for brevity the notation {/£> =/£-i/> (4.7) called the commutator of these operators. It is easily seen that {j§A = {fM+f{g,h (4.8) We notice that, if {/, h} = and {g, h) = 0, it does not in general follow that / and g commute. §5. The continuous spectrum All the relations given in §§3 and 4, describing the properties of the eigen- functions of a discrete spectrum, can be generalised without difficulty to the case of a continuous spectrum of eigenvalues. Let / be a physical quantity having a continuous spectrum. We shall 16 The Basic Concepts of Quantum Mechanics §5 denote its eigenvalues by the same letter/ simply, without suffix, in accordance with the fact that /takes a continuous range of values. We denote by T/ the eigenfunction corresponding to the eigenvalue /. Just as an arbitrary wave function Y can be expanded in a series (3.2) of eigenfunctions of a quantity having a discrete spectrum, it can also be expanded (this time as an integral) in terms of the complete set of eigenfunctions of a quantity with a continuous spectrum. This expansion has the form T( ? ) = J« / T / ( 9 )d/, (5.1) where the integration is extended over the whole range of values that can be taken by the quantity/. The subject of the normalisation of the eigenfunctions of a continuous spectrum is more complex than in the case of a discrete spectrum. The requirement that the integral of the squared modulus of the function should be equal to unity cannot here be satisfied, as we shall see below. Instead, we try to normalise the functions T/ in such a way that |«/| 2 d/is the prob- ability that the physical quantity concerned, in the state described by the wave function Y, has a value between/ and f+df. This is a direct generali- sation of the case of a discrete spectrum, where the square \a n \ 2 determines the probability of the eigenvalue f n . Since the sum of the probabilities of all possible values of / must be equal to unity, we have J>,| 2 d/=1 (5.2) (similarly to the relation (3.3) for a discrete spectrum). Proceeding in exactly the same way as in the derivation of formula (3.5), and using the same arguments, we can write, firstly, j y i n ¥*dq = j K| 2 d/ and, secondly, j YY* dq = jj a f *¥,*¥ dfdq. By comparing these two expressions we find the formula which determines the expansion coefficients, a, = JY(qy¥ f *(q)dq, (5.3) in exact analogy to (3.5). To derive the normalisation condition, we now substitute (5.1) in (5.3), and obtain a t = \a r (£¥ f W t *dq)df'. This relation must hold for arbitrary a f , and therefore must be satisfied identically. For this to be so, it is necessary that, first of all, the coefficient of a r under the integral sign (i.e. the integral J" Yj'Y/* dq) should be zero for §5 The continuous spectrum \1 all/' 7^/. For/'=/, this coefficient must become infinite (otherwise the integral over /' would vanish). Thus the integral J Yy'Y/* dq is a function of the difference /'■—/, which becomes zero for values of the argument different from zero and is infinite when the argument is zero. We denote this function by B(f'—f): JY/r,*d<z = 8(/'-/). (5.4) The manner in which the function S(/'— /) becomes infinite for/'— f= is determined by the fact that we must have J8(f'-f)a r df = a f . It is clear that, for this to be so, we must have j 8(f-f) d/' = l. The function thus defined is called a delta function, and was first used in quantum mechanics by P. A. M. Dirac. We shall write out once more the formulae which define it. They are S(«) = for x # 0, 8(0) = oo, (5.5) while f 8(x) dx = 1. (5.6) We can take as limits of integration any numbers such that x = lies between them. If f(x) is some function continuous at x = 0, then js(*)/(*)d*=/(0). (5.7) — oo * This formula can be written in the more general form j S(x-a)f(x)dx =/(«), (5.8) where the range of integration includes the point x = a, and/(#) is continuous at x = a. It is also evident that S(-x) = 8(x), (5.9) i.e. the delta function is even. Finally, writing f8(a*)d*=r8(v)^ = -, -i -i |a| |a| we can deduce that 8(«) = (1/M) S(*), (5.10) where a is any constant. The formula (5.4) gives the normalisation rule for the eigenfunctions of a continuous spectrum; it replaces the condition (3.6) for a discrete spectrum. We see that the functions T/ and Y/» with / # /' are, as before, orthogonal. 18 The Basic Concepts of Quantum Mechanics §5 However, the integrals of the squared moduli |T/| 2 of the functions diverge for a continuous spectrum. The functions Yfiq) satisfy still another relation similar to (5.4). To derive this, we substitute (5.3) in (5.1), which gives Y(ff) = J W)tf Y/WM 4/) <¥, whence we can at once deduce that we must have jV f *(q'W f (q)df=S(q'-q). (5.11) There is, of course, an analogous relation for a discrete spectrum : HY n *(q')W n (q)=*(q'-q). (5.12) n Comparing the pair of formulae (5.1), (5.4) with the pair (5.3), (5.11), we see that, on the one hand, the function x P(q) can be expanded in terms of the functions T/(^) with expansion coefficients «/ and, on the other hand, formula (5.3) represents an entirely analogous expansion of the function af = a(f) in terms of the functions x F/*(g r ), while the Y(q) play the part of expansion coefficients. The function a(f), like y ¥(q) i completely determines the state of the system; it is sometimes called a wave function in the f repre- sentation (while the function T(^) is called a wave function in the q representa- tion). Just as |T(^)| 2 determines the probability for the system to have co- ordinates lying in a given interval dq, so |«(/)| 2 determines the probability for the values of the quantity / to lie in a given interval d/. On the one hand, the functions Yfiq) are the eigenfunctions of the quantity /in the q representa- tion ; on the other hand, their complex conjugates are the eigenfunctions of the co-ordinate q in the /representation. Let<£(/) be some function of the quantity/, such that<£ and /are related in a one-to-one manner. Each of the functions Y^) can then be regarded as an eigenfunction of the quantity <f>, corresponding to a value of the latter determined by <j> = <f>(j). Here, however, the normalisation of these functions must be changed: the eigenfunctions Y^q) of the quantity <j> must be normalised by the condition jT^Wd* = 8[#/')-tf/)], whereas the functions M^, are normalised by the condition (5.4). The argu- ment of the delta function becomes zero only for/' = /. As /' approaches/, we have cf>(f') -<f>(f) = [d<£(/)/d/] . (/' -/). By (5.10) we can therefore writef «/')-<£(/)] = [AM( \ IA J U'-f)- ( 5 -!3) t In general, if<f>(x) is some one- valued function (the inverse function need not be one-valued), we have 8[<f>(x)] = V S(x— aA where a,- are the roots of the equation <f>(x) = 0. §5 The continuous spectrum 19 Thus the normalisation condition for the functions T^ can be written in the form l *v>Y„f d q = — -^— -scr-/). l<W)/d/l Comparing this with (5.4), we see that the functions T^ and ¥> are related by r * w =vmw? f - (5 - 14) There are also physical quantities which in one range of values have a discrete spectrum, and in another a continuous spectrum. For the eigen- functions of such a quantity all the relations derived in this and the previous sections are, of course, true. It need only be noted that the complete set of functions is formed by combining the eigenfunctions of both spectra. Hence the expansion of an arbitrary wave function in terms of the eigenfunc- tions of such a quantity has the form Yfe) = S a n r n (q)+ j a f Y f (q) d/, (5.15) where the sum is taken over the discrete spectrum and the integral over the whole continuous spectrum. The co-ordinate q itself is an example of a quantity having a continuous spectrum. It is easy to see that the operator corresponding to it is simply multiplication by q. For, since the probability of the various values of the co-ordinate is determined by the square |Y(?)| 2 , the mean value of the co-ordinate is q = J q\W\ 2 dq. On the other hand, the mean value of the co-ordinate must be determined by its operator as q^^T^qYdq. A comparison of the two expressions shows that the operator q is simply multiplication by q\ this may be symbolically written in the formf § = q- (5.16) The eigenfunctions of this operator must be determined, according to the usual rule, by the equation qT g<i = q Y qo , where q temporarily denotes the actual values of the co-ordinate as distinct from the variable q. Since this equation can be satisfied either by Y q<i = or by q = q , it is clear that the eigenfunctions which satisfy the normalisation condition arej % o = S(q-q ). (5.17) t In future we shall always, for simplicity, write operators which amount to multiplication by some quantity in the form of that quantity itself. J The expansion coefficients for an arbitrary function Y in terms of these eigenfunctions are \ = \nm<i-qo)dq = V(qo). The probability that the value of the co-ordinate lies in a given interval dq is W<, \ 2 dq = mq )\ z dq , as it should be. 20 The Basic Concepts of Quantum Mechanics §6 §6. The passage to the limiting case of classical mechanics Quantum mechanics contains classical mechanics in the form of a certain limiting case. The question arises as to how this passage to the limit is made. In quantum mechanics an electron is described by a wave function which determines the various values of its co-ordinates ; of this function we so far know only that it is the solution of a certain linear partial differential equation. In classical mechanics, on the other hand, an electron is regarded as a material particle, moving in a path which is completely determined by the equations of motion. There is an interrelation, somewhat similar to that between quantum and classical mechanics, in electrodynamics between wave optics and geometrical optics. In wave optics, the electromagnetic waves are described by the electric and magnetic field vectors, which satisfy a definite system of linear differential equations, namely Maxwell's equations. In geometrical optics, however, the propagation of light along definite paths, or rays, is considered. Such an analogy enables us to see that the passage from quantum mechanics to the limit of classical mechanics occurs similarly to the passage from wave optics to geometrical optics. Let us recall how this latter transition is made mathematically. Let u be any of the field components in the electromagnetic wave. It can be written in the form u = ae i( f> (with a and<£ real), where a is called the amplitude and<£ the phase of the wave. The limiting case of geometrical optics corresponds to small wavelengths; this is expressed mathematically by saying that <f> (called in geometrical optics the eikonal) varies by a large amount over short distances ; this means, in particular, that it can be supposed large in absolute value. Similarly, we start from the hypothesis that, to the limiting case of classical mechanics, there correspond in quantum mechanics wave functions of the form T = ae**!*, where a is a slowly varying function and <j> takes large values. As is well known, the path of a particle can be determined in mechanics by means of the variational principle, according to which what is called the action S of a mechanical system must take its least possible value (the principle of least action, or Hamilton's principle). In geometrical optics the path of the rays is determined by what is called Fermat's principle, according to which the optical path length of the ray, i.e. the difference between its phases at the beginning and end of the path, must take its least (or greatest) possible value. On the basis of this analogy, we can assert that the phase <j> of the wave function, in the limiting (classical) case, must be proportional to the mech- anical action S of the physical system considered, i.e. we must have S = constant X <f>. The constant of proportionality is called Planck's con- stanff and is denoted by h. It has the dimensions of action (since <f> is t It was introduced into physics by M. Planck in 1900. The constant h, which we use everywhere in this book, is, strictly speaking, Planck's constant divided by 2n. §7 The wave function and measurements 21 dimensionless) and has the value h = 1-054 xlO- 27 erg sec. Thus, the wave function of an "almost classical" (or, as we say, quasi- classical) physical system has the form T = ae^/K. (6.1) Planck's constant h plays a fundamental part in all quantum phenomena. Its relative value (compared with other quantities of the same dimensions) determines the "extent of quantisation" of a given physical system. The transition from quantum mechanics to classical mechanics, corresponding to large phase, can be formally described as a passage to the limit h ->0 (just as the transition from wave optics to geometrical optics corresponds to a passage to the limit of zero wavelength, A -> 0). We have ascertained the limiting form of the wave function, but the question still remains how it is related to classical motion in a path. In general, the motion described by the wave function does not tend to motion in a definite path. Its connection with classical motion is that, if at some initial instant the wave function, and with it the probability distribution of the co-ordinates, is given, then at subsequent instants this distribution will change according to the laws of classical mechanics (for a more detailed dis- cussion of this, see the end of §17). In order to obtain motion in a definite path, we must start from a wave function of a particular form, which is perceptibly different from zero only in a very small region of space (what is called a wave packet); the dimensions of this region must tend to zero with h. Then we can say that, in the quasi- classical case, the wave packet will move in space along a classical path of a particle. Finally, quantum-mechanical operators must reduce, in the limit, simply to multiplication by the corresponding physical quantity. §7. The wave function and measurements Let us again return to the process of measurement, whose properties have been qualitatively discussed in §1 ; we shall show how these properties are related to the mathematical formalism of quantum mechanics. We consider a system consisting of two parts: a classical apparatus and an electron (regarded as a quantum object). The process of measurement consists in these two parts' coming into interaction with each other, as a result of which the apparatus passes from its initial state into some other; from this change of state we draw conclusions concerning the state of the' electron. The states of the apparatus are distinguished by the values of some physical quantity (or quantities) characterising it— the "readings of the ap- paratus". We conventionally denote this quantity by g, and its eigenvalues by g n ; these take in general, in accordance with the classical nature of the apparatus, a continuous range of values, but we shall— merely in order to 22 The Basic Concepts of Quantum Mechanics §7 simplify the subsequent formulae — suppose the spectrum discrete. The states of the apparatus are described by means of quasi-classical wave func- tions, which we shall denote by <& n (£), where the suffix n corresponds to the "reading" g n of the apparatus, and £ denotes the set of its co-ordinates. The classical nature of the apparatus appears in the fact that, at any given instant, we can say with certainty that it is in one of the known states <D n with some definite value of the quantity g\ for a quantum system such an assertion would, of course, be unjustified. Let O (£) be the wave function of the initial state of the apparatus (before the measurement), and T(^) some arbitrary normalised initial wave function of the electron (q denoting its co-ordinates). These functions describe the state of the apparatus and of the electron independently, and therefore the initial wave function of the whole system is the product Y(?)O (|). (7.1) Next, the apparatus and the electron interact with each other. Applying the equations of quantum mechanics, we can in principle follow the change of the wave function of the system with time. After the measuring process it may not, of course, be a product of functions of £ and q. Expanding the wave function in terms of the eigenfunctions O n of the apparatus (which form a complete set of functions), we obtain a sum of the form XA n {q)® n (Z), (7-2) where the A n (q) are some functions of q. The classical nature of the apparatus, and the double role of classical mechanics as both the limiting case and the foundation of quantum mechanics, now make their appearance. As has been said above, the classical nature of the apparatus means that, at any instant, the quantity g (the "reading of the apparatus") has some definite value. This enables us to say that the state of the system apparatus + electron after the measurement will in actual fact be described, not by the entire sum (7.2), but by only the one term which corresponds to the "reading" g n of the apparatus, 4.(«)*.tf). ( 7 - 3 ) It follows from this that A n {q) is proportional to the wave function of the electron after the measurement. It is not the wave function itself, as is seen from the fact that the function A n (q) is not normalised. It contains both information concerning the properties of the resulting state of the electron and the probability (determined by the initial state of the system) of the occurrence of the nth "reading" of the apparatus. Since the equations of quantum mechanics are linear, the relation between A n {q) and the initial wave function of the electron Y(q) is in general given by some linear integral operator : A n (q) = JK n (q,q'W(q')dq', (7.4) §7 The wave function and measurements 23 with a kernel K n (q, q') which characterises the measurement process con- cerned. We shall suppose that the measurement concerned is such that it gives a complete description of the state of the electron. In other words (see §1), in the resulting state the probabilities of all the quantities must be indepen- dent of the previous state of the electron (before the measurement). Mathe- matically, this means that the form of the functions A n (q) must be determined by the measuring process itself, and does not depend on the initial wave function Y(q) of the electron. Thus the A n must have the form Mfi = " n <f>n(q), (7.5) where the <f> n are definite functions, which we suppose normalised, and only the constants a^ depend on T(^). In the integral relation (7.4) this corresponds to a kernel K n {q, q') which is a product of a function of q and a function of q' : Kn{q,q') = <i>n{qWn*(q'y, (7.6) then the linear relation between the constants a n and the function Y(q) is a^J^WF.-fo)^, (7.7) where the T n (^) are certain functions depending on the process of measure- ment. The functions <j> n {q) are the normalised wave functions of the electron after measurement. Thus we see how the mathematical formalism of the theory reflects the possibility of finding by measurement a state of the electron described by a definite wave function. If the measurement is made on an electron with a given wave function f (q), the constants a n have a simple physical meaning: in accordance with the usual rules, |aj 2 is the probability that the measurement will give the nth result. The sum of the probabilities of all results is equal to unity: ?KI 2 = 1. (7.8) In order that equations (7.7) and (7. 8) should hold for an arbitrary normalised function W(q\ it is necessary (cf. §3) that an arbitrary function T(g) can be expanded in terms of the functions Y n (y). This means that the functions ^n(q) form a complete set of normalised and orthogonal functions. If the initial wave function of the electron coincides with one of the func- tions Y n (g), then the corresponding constant a n is evidently equal to unity, while all the others are zero. In other words, a measurement made on an electron in the state T n (#) gives with certainty the nth result. All these properties of the functions Y n (q) show that they are the eigen- functions of some physical quantity (denoted by /) which characterises the electron, and the measurement concerned can be spoken of as a measurement of this quantity. It is very important to notice that the functions T TC (^) do not, in general, 24 The Basic Concepts of Quantum Mechanics §7 coincide with the functions <f> n (q) ; the latter are in general not even mutually orthogonal, and do not form a set of eigenfunctions of any operator. This expresses the fact that the results of measurements in quantum mechanics cannot be reproduced. If the electron was in a state T n (^), then a measure- ment of the quantity / carried out on it leads with certainty to the value f n . After the measurement, however, the electron is in a state c/) n (q) different from its initial one, and in this state the quantity / does not in general take any definite value. Hence, on carrying out a second measurement on the electron immediately after the first, we should obtain for /a value which did not agree with that obtained from the first measurement.-f To predict (in the sense of calculating probabilities) the result of the second measurement from the known result of the first, we must take from the first measurement the wave function <j> n {q) of the state in which it resulted, and from the second measurement the wave function T n (^) of the state whose probability is re- quired. This means that from the equations of quantum mechanics we deter- mine the wave function c/> n (q, t) which, at the instant when the first measure- ment is made, is equal to<£ n (<?) ; the probability of the mth result of the second measurement, made at time t, is then given by the squared modulus of the integral J^ n ( ?l t)Y m *(q) 6q. We see that the measuring process in quantum mechanics has a "two- faced" character: it plays different parts with respect to the past and future of the electron. With respect to the past, it "verifies" the probabilities of the various possible results predicted from the state brought about by the previ- ous measurement. With respect to the future, it brings about a new state (see also §44). Thus the very nature of the process of measurement involves a far-reaching principle of irreversibility. This irreversibility is of fundamental significance. We shall see later (at the end of §18) that the basic equations of quantum mechanics are in them- selves symmetrical with respect to a change in the sign of the time; here quantum mechanics does not differ from classical mechanics. The irrever- sibility of the process of measurement, however, causes the two directions of time to be physically non-equivalent, i.e. creates a difference between the future and the past. t It must be remarked that there is an important exception to the statement that results of measure- ments cannot be reproduced : the one quantity the result of whose measurement can be exactly re- produced is the co-ordinate. Two measurements of the co-ordinates of an electron, made at a sufficiently small interval of time, must give neighbouring values; if this were not so, it would mean that the electron had an infinite velocity. Mathematically, this is related to the fact that the co-ordinate commutes with the operator of the interaction energy between the electron and the apparatus, since this energy is (in non-relativistic theory) a function of the co-ordinates only. CHAPTER II ENERGY AND MOMENTUM §8. The Hamiltonian operator The wave function Y completely determines the state of a physical system in quantum mechanics. This means that, if this function is given at some instant, not only are all the properties of the system at that instant described, but its behaviour at all subsequent instants is determined (only, of course, to the degree of completeness which is generally admissible in quantum mech- anics). The mathematical expression of this fact is that the value of the deri- vative <W/dt of the wave function with respect to time at any given instant must be determined by the value of the function itself at that instant, and, by the principle of superposition, the relation between them must be linear. In the most general form we can write where L is some linear operator; the factor i is introduced here for conveni- ence. We shall derive some properties of the operator L. Since the integral J YY* dq is a constant independent of time, we have Substituting here dYjdt = -iD¥, dY*{dt = it*W* and using in the first integral the definition of the transpose of an operator, we can write J YZ*Y* dq - J Y*&F dq = j Y*Z*Y dq - j >¥*&¥ dq = J x F*(l*-l)Wdq = 0. Since this equation must hold for an arbitrary function T, it follows that we must have identically Z* — L = 0, or The operator L is therefore Hermitian. Let us find the classical quantity to which the operator L corresponds. To do this, we use the limiting expression (6.1) for the wave function and write BY _ i dS ~dt~h~dt ' the slowly varying amplitude a need not be differentiated. Comparing this 25 26 Energy and Momentum §9 equation with the definition dYjdt = —iD¥, we see that, in the limiting case, the operator L reduces to simply multiplying by — (l/#) dSjdt. This means that —(1//*) dSjdt is the physical quantity into which the Hermitian operator L passes. As is well known from mechanics, the derivative —dSjdt is just Hamilton's function H for a mechanical system. Thus the operator hL is what corres- ponds in quantum mechanics to Hamilton's function; this operator, which we shall denote by i?, is called the Hamiltonian operator or, more briefly, the Hamiltonian of the system. The relation between dY/dt and T is ihd y ¥ldt=fi K ¥. (8.1) If the form of the Hamiltonian is known, equation (8.1) determines the wave functions of the physical system concerned. This fundamental equation of quantum mechanics is called the wave equation. §9. The differentiation of operators with respect to time The concept of the derivative of a physical quantity with respect to time cannot be denned in quantum mechanics in the same way as in classical mech- anics. For the definition of the derivative in classical mechanics involves the consideration of the values of the quantity at two neighbouring but distinct instants of time. In quantum mechanics, however, a quantity which at some instant has a definite value does not in general have definite values at subsequent instants; this was discussed in detail in §1. Hence the idea of the derivative with respect to time must be differently defined in quantum mechanics. It is natural to define the derivative / of a quantity / as the quantity whose mean value is equal to the derivative, with respect to time, of the mean value /. Thus we have the definition /=/ (9-1) Starting from this definition, it is easy to obtain an expression for the quantum-mechanical operator / corresponding to the quantity /. Since f = jY*fYdq, Here dfjdt is the operator obtained by differentiating the operator / with respect to time; /may depend on the time as a parameter. Substituting for BYjdt, dW^/dt their expressions according to (8.1), we obtain /= [V*— Wdq+- f (#*Y*)/Td?-^ (wf(fr¥)dq. J dt hj n J Since the operator i? is Hermitian, we have f (i?*T*)(/T) dq = JT*^/T dq: §10 Stationary states 27 thus / -J**S + i* L >> d »- Since, on the other hand, we must have, by the definition of mean values, / = J Y*/ Tdg-, it is seen that the expression in parentheses under the integral is the required operator /:f We notice that, if the operator / is independent of time, / reduces, apart from a constant factor, to the result of commuting the operator / with the Hamiltonian. A very important class of physical quantities is formed by those whose operators do not depend explicitly on time, and also commute with the Hamiltonian, so that / = 0. Such quantities are said to be conserved. If the operator /is identically zero, then/ =/= 0, that is, /is constant. In other words, the mean value of the quantity remains constant in time. We can also assert that, if in a given state the quantity /has a definite value (i.e. the wave function is an eigenfunction of the operator /), then it will have a definite value (the same one) at subsequent instants also. §10. Stationary states If the system is not in a varying external field, its Hamiltonian cannot contain the time explicitly. This follows at once from the fact that, in the :+ 5& + |f<> f In classical mechanics we have for the total derivative, with respect to time, of a quantity / which is a function of the generalised co-ordinates g, and momenta p { of the system dt dt Substituting, in accordance with Hamilton's equations, q t = dH/dpi and p t = —dHldq t , we obtain d//d* = df/dt+[H,f], where Z-w \dq t dpi dpi dqtJ i is what is called the Poisson bracket for the quantities /and H (see Mechanics, §42). On comparing with the expression (9.2), we see that, as we pass to the limit of classical mechanics, the operator i($f—fff) reduces in the first approximation to zero, as it should, and in the second approximation (with respect to K) to the quantity H[H,f]. This result is true also for any two quantities / and g; the operator i(fg-gf) tends in the limit to the quantity h[f, g] , where [/, g] is the Poisson bracket ^L,\dqt dpi dpidqt)' This follows at once from the fact that we can always formally imagine a system whose Hamiltonian 28 Energy and Momentum §10 absence of an external field (or in a constant external field) all times are equivalent so far as the given physical system is concerned. Since, on the other hand, any operator of course commutes with itself, we reach the con- clusion that Hamilton's function is conserved for systems which are not in a varying external field. As is well known, a Hamilton's function which is conserved is called the energy. Thus we have the law of conservation of energy in quantum mechanics. Here it signifies that, if in a given state the energy has a definite value, this value remains constant in time. States in which the energy has definite values are called stationary states of a system. They are described by wave functions T n which are the eigen- functions of the Hamiltonian operator, i.e. which satisfy the equation i?T n = E^¥ n , where E n are the eigenvalues of the energy. Correspondingly, the wave equation (8.1) for the function T n , ihdW n ldt=8Y n = E n y n can be integrated at once with respect to time and gives T n = erU/*>*:ntf n (q), (10.1) where ip n is a function of the co-ordinates only. This determines the relation between the wave functions of stationary states and the time. We shall denote by the small letter i/j the wave functions of stationary states without the time factor. These functions, and also the eigenvalues of the energy, are determined by the equation ify = E+. (10.2) The stationary state with the smallest possible value of the energy is called the normal or ground state of the system. The expansion of an arbitrary wave function \F in terms of the wave func- tions of stationary states has the form T = S a n e-QM E «tif,Jq). (10.3) n The squared moduli \a n \ 2 of the expansion coefficients, as usual, determine the probabilities of various values of the energy of the system. The probability distribution for the co-ordinates in a stationary state is determined by the squared modulus pFJ 2 = |</r n | 2 ; we see that it is indepen- dent of time. The same is true of the mean values /=jT n */Y n d ? = J^ n */0 n d 9 of any physical quantity / (whose operator does not depend explicitly on the time), and therefore of the probabilities of its various values. As has been said, the operator of any quantity that is conserved commutes with the Hamiltonian. This means that any physical quantity that is con- served can be measured simultaneously with the energy. §10 Stationary states 29 Among the various stationary states, there may be some which correspond to the same value of the energy, but differ in the values of some other physical quantities. Such eigenvalues of the energy (or, as we say, energy levels of the system), to which several different stationary states correspond, are said to be degenerate. Physically, the possibility that degenerate levels can exist is related to the fact that the energy does not in general form by itself a complete set of physical quantities. In particular, it is easy to see that, if there are two conserved physical quantities /and g whose operators do not commute, then the energy levels of the system are in general degenerate. For, let $ be the wave function of a stationary state in which, besides the energy, the quantity /also has a definite value. Then we can say that the function gifi does not coincide (apart from a constant factor) with ip; if it did, this would mean that the quantity g also had a definite value, which is impossible, since / and g cannot be measured simultaneously. On the other hand, the function gip is an eigenfunction of the Hamiltonian, corresponding to the same value E of the energy as ifs: &m =§&+ = %&!>)• Thus we see that the energy E corresponds to more than one eigenfunction, i.e. the energy level is degenerate. It is clear that any linear combination of wave functions corresponding to the same degenerate energy level is also an eigenfunction for that value of the energy. In other words, the choice of eigenfunctions of a degenerate energy level is not unique. Arbitrarily selected eigenfunctions of a degener- ate energy level are not, in general, orthogonal. By a proper choice of linear combinations of them, however, we can always obtain a set of orthogonal (and normalised) eigenfunctions (and this can be done in infinitely many ways; for the number of independent coefficients in a linear transformation of n functions is n 2 , while the number of normalisation and orthogonality conditions for n functions is \n{n-\-\), i.e. less than w 2 ). These statements concerning the eigenfunctions of a degenerate energy level relate, of course, not only to eigenfunctions of the energy, but also to those of any operator. Thus only those functions are automatically ortho- gonal which correspond to different eigenvalues of the operator concerned ; functions which correspond to the same degenerate eigenvalue are not in general orthogonal. If the Hamiltonian of the system is the sum of two (or more) parts, H = ^+^2, one of which contains only the co-ordinates q x and the other only the co-ordinates q 2 , then the eigenfunctions of the operator H can be written down as products of the eigenfunctions of the operators 3 X and i? 2 , and the eigenvalues of the energy are equal to the sums of the eigenvalues of these operators. The spectrum of eigenvalues of the energy may be either discrete or continuous. A stationary state of a discrete spectrum always corresponds to a finite motion of the system, i.e. one in which neither the system nor any 30 Energy and Momentum §11 part of it moves off to infinity. For, with eigenfunctions of a discrete spec- trum, the integral j |Y| 2 dq, taken over all space, is finite. This certainly means that the squared modulus |Y| 2 decreases quite rapidly, becoming zero at infinity. In other words, the probability of infinite values of the co- ordinates is zero ; that is, the system executes a finite motion, and is said to be in a bound state. For wave functions of a continuous spectrum, the integral J" |T| 2 dq diverges. Here the squared modulus |T| 2 of the wave function does not directly deter- mine the probability of the various values of the co-ordinates, and must be regarded only as a quantity proportional to this probability. The divergence of the integral J" |T| 2 dq is always due to the fact that |T| 2 does not become zero at infinity (or becomes zero insufficiently rapidly). Hence we can say that the integral J" |*F| 2 dq, taken over the region of space outside any arbi- trarily large but finite closed surface, will always diverge. This means that, in the state considered, the system (or some part of it) is at infinity. For a wave function which is a superposition of the wave functions of various stationary states of a continuous spectrum, the integral J" pF| 2 dq may converge, so that the system lies in a finite region of space. However, it can be shown that, in the course of time, this region moves unrestrictedly, and eventually, the system moves off to infinity.^ Thus the stationary states of a continuous spectrum correspond to an infinite motion of the system. §11. Matrices We shall suppose for convenience that the system considered has a discrete energy spectrum ; all the relations obtained below can be generalised at once to the case of a continuous spectrum. Let *P = 1,a n , ¥ n be the expansion of an arbitrary wave function in terms of the wave functions \F n of the stationary states. If we substitute this expansion in the definition /= jW^f^V dq of | This can be seen as follows. The superposition of wave functions of a continuous spectrum has the form T = f aEe-WEtyEiq) d£. The squared modulus of *F can be written in the form of a double integral: |T| 2 = J* J" a E a E '*^l^ E '- E ^iljE{q)4'E-*{q) dEdE'. If we average this expression over some time interval T, and then let T tend to infinity, the mean values of the oscillating factors eC/AK- 8 '-*) 1 , and therefore the whole integral, tend to zero in the limit. Thus the mean value, with respect to time, of the probability of finding the system at any given point of configuration space tends to zero. This is possible only if the motion takes place throughout infinite space. We note that, for a function *F which is a superposition of functions of a discrete spectrum, we should have |T| 2 = SS a n a m *e^l^ m -^t^ m * = S \a n ^ n {q)\\ nm n i.e. the required probability remains finite on averaging over time. §11 Matrices 31 the mean value of some quantity /, we obtain f=maSa m f um (t)> (H.1) n m where f nm (t) denotes the integral fnJt) = \V n *fV m dq. (11.2) The set of quantities f nm (t) with all possible n and m is called the matrix^ of the quantity/, and each of the/ nm (i) is called the matrix element corresponding to the transition from state n to state m.% The dependence of the matrix elements f nm (f) on time is determined (if the operator / does not contain the time explicitly) by the dependence of the functions T n on time. Substituting for them the expressions (10.1), we find that UJt) =f nm e ia, ™ t , (11.3) where "nm = (E n -E m )/h (11.4) is what is called the transition frequency between the states n and m, and the quantities f nm = j<l>«*fK*<l (H.5) form the matrix of the quantity / which is independent of time, and which is commonly used. j| We note that the "frequencies" co nm satisfy the ob- vious relation w» m +w«i = w„|. (11.6) The matrix elements of the derivative / are obtained by differentiating the matrix elements of the quantity / with respect to time; this follows directly from the fact that the mean value / is equal to /, i.e. /= 22 a n *a m f ntn (t). nm From (11.3) we thus have for the matrix elements of/ fnmit) = i<*>nmfn m (t) (H-7) or (cancelling the time factor e <a W from both sides) for the matrix elements independent of time (/)«m = io>nmfnm = (ijH)(E n -E m )f nm . (11,8) f The matrix representation of physical quantities was introduced by Heisenberg in 1925, before Schrodinger's discovery of the wave equation. "Matrix mechanics" was later developed by M. Born, W. Heisenberg and P. Jordan. J In some cases, when each of the suffixes n and m has to be written in the form of several letters, we shall use the notation / m n instead of/ nm . The notation (n\f \m) is also used. || It must be borne in mind that, because of the indeterminacy of the phase factor in normalized Wave functions (see §2), the matrix elements/„ m (and/ nflt (i)) also are determined only to within a factor of the form eihm-ar). Here again this indeterminacy has no effect on any physical results. 32 Energy and Momentum §11 To simplify the notation in the formulae, we shall derive all our relations below for the matrix elements independent of time ; exactly similar relations hold for the matrices which depend on the time. For the matrix elements of the complex conjugate/* of the quantity/ we obtain, taking into account the definition of the Hermitian conjugate operator, (/•).« = / 4>n*f + *l> m dq = j + n *f*K dq = j tJ*t n * dq - or (f*)nm = (fmn)*. (11.9) For real physical quantities, which are the only ones we usually consider, we consequently have /nra = /m« (11.10) {fmn* stands for (f mn )*). Such matrices, like the corresponding operators, are said to be Hermitian. Matrix elements with n = m are called diagonal elements. These are independent of time, and (11.10) shows that they are real. The element f nn is the mean value of the quantity /in the state ip n . It is not difficult to obtain the "multiplication rule" for matrices. To do so, we first observe that the formula fy n = $f mn if> m (11.11) holds. This is simply the expansion of the function fip n in terms of the func- tions ift m , the coefficients being determined in accordance with the general formula (3.5). Remembering this formula, let us write down the result of the product of two operators acting on the function ip n : fg'Pn = ftg*pn) = / S g kn *fr k = 2 g kn fy k =2 g kn f mk ^ m . ic k fc,m Since, on the other hand, we must have fgtn = S {fg) m rrtm, we arrive at the result that the matrix elements of the product fg are deter- mined by the formula (fg)mn = Xfmkgkn- (H-12) This rule is the same as that used in mathematics for the multiplication of matrices. If the matrix is given, then so is the operator itself. In particular, if the matrix is given, it is in principle possible to determine the eigenvalues of the physical quantity concerned and the corresponding eigenf unctions. We shall now consider the values of all quantities at some definite instant, and expand an arbitrary wave function \F (at that instant) in terms of the eigenfunctions of Hamilton's operator H, i.e. of the wave functions ip m of the stationary states (these wave functions are independent of time) : T=S^ m , (11.13) m §11 Matrices 33 where the expansion coefficients are denoted by c m . We substitute this expan- sion in the equation /T =f¥ which determines the eigenvalues and eigen- functions of the quantity/. We have 2 Cmifym) —f 2 C m ijj m . m m We multiply both sides of this equation by iff n * and integrate over q. Each of the integrals J" i/* n *fy m d<7 on the left-hand side of the equation is the cor- responding matrix element f nm . On the right-hand side, all the integrals J il/ n *tp m &q with m ^ n vanish by virtue of the orthogonality of the functions tp m , and J n *j/f n dq = 1 by virtue of their normalisation. Thus ^JnmCm == J^m m or S(/„ m -/S nm V m = 0, (11.14) m where 8 nm = for m # n and = 1 for m = n. Thus we have obtained a system of homogeneous algebraic equations of the first degree (with the c m as unknowns). As is well known, such a system has solutions different from zero only if the determinant formed by the coefficients in the equations vanishes, i.e. only if I fnm—finml = 0- The roots of this equation (in which / is regarded as the unknown) are the possible values of the quantity/. The set of values c m satisfying the equations (11.14) when /is equal to any of these values determines the corresponding eigenf unction. If, in the definition (11.5) of the matrix elements of the quantity/, we take as ift n the eigenfunctions of this quantity, then from the equation ftfi n =f n ijj n we have fnm = J «An*/«A«» d? = f m J >f> n *tp m d?. By virtue of the orthogonality and normalisation of the functions j/r m , this gives f nm = for n ^ m and f mm = f m . Thus only the diagonal matrix elements are different from zero, and each of these is equal to the correspond- ing eigenvalue of the quantity/. A matrix with only these elements different from zero is said to be put in diagonal form. In particular, in the usual representation, with the wave functions of the stationary states as the functions iff n , the energy matrix is diagonal (and so are the matrices of all other physical quantities having definite values in the stationary states). In general, the matrix of a quantity /, defined with respect to the eigenfunctions of some operator £, is said to be the matrix of fin a representation in which g is diagonal. We shall always, except where the subject is specially mentioned, understand in future by the matrix of a physical quantity its matrix in the usual repre- sentation, in which the energy is diagonal. Everything that has been said 34 Energy and Momentum §11 above regarding the dependence of matrices on time refers, of course, only to this usual representation.f By means of the matrix representation of operators we can prove the theorem mentioned in §4: if two operators commute with each other, they have their entire sets of eigenfunctions in common. Let / and g be two such operators. From fg = gf and the matrix multiplication rule (11.12), it follows that jfjmkSkn = 4? gmkjkn- If we take the eigenfunctions of the operator /as the set of functions ift n with respect to which the matrix elements are calculated, we shall have f mk = for m # k, so that the above equation reduces to f mm g mn = g mn f nnl or SmnKJm Jn) == U. If all the eigenvalues f n of the quantity / are different, then for all m ^n we nave fm—fn # 0» so that we must have g mn = 0. Thus the matrix g mn is also diagonal, i.e. the functions ifj n are eigenfunctions of the physical quantity g also. If, among the values / n , there are some which are equal (i.e. if there are eigenvalues to which several different eigenfunctions correspond), then the matrix elements g mn corresponding to each such group of functions tp n are, in general, different from zero. However, linear combinations of the functions ifi n which correspond to a single eigenvalue of the quantity / are evidently also eigenfunctions of/; one can always choose these combinations in such a way that the corresponding non-diagonal matrix elements g mn are zero, and thus, in this case also, we obtain a set of functions which are simultaneously the eigenfunctions of the operators / and g. The matrix f nm can be regarded as the operator /in the energy representa- tion. For the set of coefficients c n in the expansion (11.13) in terms of the eigenfunctions ip n of the Hamiltonian can be considered (cf. §5) as the wave function in the "E representation" (the variable being the suffix n which gives the number of the stationary state). The formula n m for the mean value of the quantity/ then corresponds to the general expression for the quantum-mechanical mean value of a quantity in terms of its operator and the wave function of the state concerned. PROBLEM The Hamiltonian of a system, and therefore the eigenvalues E n of the energy, are functions of some parameter A. Show that (0H/SA)„„ = dE n ldX. Solution. Differentiating the equation (&—E n )>]>n = with respect to A and then f Bearing in mind the diagonality of the energy matrix, it is easy to see that equation (11.8) is the operator relation (9.2) written in matrix form. §12 Transformation of matrices 35 multiplying on the left by <fi n *, we obtain On integration with respect to q, the left-hand side gives zero, since the operator i? is Hermitian and therefore J rf, n *(fi-E n )—dq = J ~{n-E n )*4, n *dq. The right-hand side gives the required equation. §12. Transformation of matrices The matrix elements of a given physical quantity can be defined with respect to various sets of wave functions, for example the wave functions of stationary states described by various sets of physical quantities, or the wave functions of stationary states of the same system in various external fields. The problem therefore arises of the transformation of matrices from one representation to another. Let ip n (q) and ifs n '{q) (n = 1,2,...) be two complete sets of orthonormal functions, related by some linear transformation : ifjn = 2 S mn tfi m , (12.1) which is simply an expansion of the function i// n ' in terms of the complete set of functions ip n . This transformation may be conventionally written in the operator form «£«' = fyn. (12.2) The operator S must satisfy a certain condition in order that the functions tpn should be orthonormal if the functions ifs n are. Substituting (12.2) in the condition J tym*$n &q = 8 m n, and using the definition of the transposed operator (3.14), we have j (Sfa)£f m * dq = j $„*§•£$* dq = B mn . If these equations hold for all m and n, we must have §*§ = 1, or £* = §+ = £-i f (12.3) i.e. the inverse operator is equal to the Hermitian conjugate operator. Operators having this property are said to be unitary. Owing to this property, the transformation iji n = S- 1 ^' inverse to (12.1) is given by tffn = 2 S nm *xfj m '. (12.4) 36 Energy and Momentum §12 Writing the equations 3 + S = 1 and SS + = 1 in matrix form, we obtain the following forms of the unitarity condition : S Sim* Sin = 8 mn , (12.5) E S m i*S n i = 8 mn . (12.6) Let us now consider some physical quantity / and write down its matrix elements in the "new" representation, i.e. with respect to the functions ip n '. These are given by the integrals j 0m'*$n' d? = J (S*iP m *)(fSiP n ) dq = J ^m*S*f§iJj n dq = jif>m*S-lfSifj n dq. Hence we see that the matrix of the operator / in the new representation is equal to the matrix of the operator /' = S-ifS (12.7) in the old representation. The sums of the diagonal elements of matrices are of importance in certain calculations in quantum mechanics. Such a sum is called the trace or spur\ of the matrix and denoted by tr/: tr/=2/ ww . (12.8) It may be noted first of all that the trace of a product of two matrices is independent of the order of multiplication : trtfc) = trfe/), (12-9) since the rule of matrix multiplication gives tr (fg) = 2 Yffnkgkn = S "Zg/cnfnk = tr (gf). Similarly we can easily see that, for a product of several matrices, the trace is unaffected by a cyclic permutation of the factors; for example, tr(fgh) = tr(hfg) = tr(ghf). (12.10) An important property of the trace is that it does not depend on the choice of the set of functions with respect to which the matrix elements are defined, since (tr/)' = tr (S-ifS) = tr (SS^f) = tr/. t From the German word Spur. The notation sp / is also used. The trace can be defined, of course, only if the sum over n is convergent; we shall assume that this condition is satisfied. §13 The Heisenberg representation of operators 37 §13. The Heisenberg representation of operators In the mathematical formalism of quantum mechanics described here, the operators corresponding to various physical quantities act on functions of the co-ordinates and do not usually depend explicitly on time. The time depen- dence of the mean values of physical quantities is due only to the time dependence of the wave function of the state, according to the formula f(t) = jW*(q,t)/T(q,t)dq. (13.1) The quantum-mechanical treatment can, however, be formulated also in a somewhat different but equivalent form, in which the time dependence is transferred from the wave functions to the operators. Although we shall not use this Heisenberg representation of operators in the present volume, a state- ment of it is given here with a view to applications in the relativistic theory. We define the operator S = exp[-(*/£)#*], ( 13-2 ) where R is the Hamiltonian of the system. By definition, its eigenfunctions are the same as those of the operator fi, i.e. the stationary-state wave functions *ffn(q)y where %(<?) = e-wn)E n t Mq) , (13>3) Hence it follows that the expansion (10.3) of an arbitrary wave function T in terms of the stationary-state wave functions can be written in the operator form V(q,t) = &¥(q,0), (13.4) i.e. the effect of the operator £ is to convert the wave function of the system at some initial instant into the wave function at an arbitrary instant. According to the definition (3.16), using the fact that the operator fi is Hermitian, we have S+ = ex pfeM = expZ-tfA = £-1, i.e. 3 is a unitary operator, as it should be, since formula (13.4) (with the time t as a parameter) may be regarded as a particular case of the transforma- tion (12.1). Defining, as in (12.7), the time-dependent operator , /(*) = &-W, (13.5) we have v ' f(t) = jV*(q,0)f(t)Y(q,0)dq, (13. 6 ) and thus obtain the formula (3.8) for the mean value of the quantity / in a form in which the time dependence is entirely transferred to the operator (for our definition of an operator rests on formula (3.8)). 38 Energy and Momentum §14 It is evident that the matrix elements of the operator (13.5) with respect to the stationary-state wave functions are the same as the time-dependent matrix elements fnm(t) defined by formula (11.3). Finally, differentiating the expression (13.5) with respect to time (assuming that the operators / and H do not themselves involve t), we obtain U{t) = Utif-fti), (13.7) ot n which is similar in form to (9.2) but has a somewhat different significance: the expression (9.2) defines the operator / corresponding to the physical quantity/, while the left-hand side of equation (13.7) is the time derivative of the operator of the quantity / itself. §14. The density matrix Let us consider a system which is a part of some closed system. We suppose that the closed system as a whole is in some state described by the wave function Y(<7, x), where x denotes the set of co-ordinates of the system considered, and q the remaining co-ordinates of the closed system. This function in general does not fall into a product of functions of x and of q alone, so that the system does not have its own wave function, f Let / be some physical quantity pertaining to the system considered. Its operator therefore acts only on the co-ordinates x, and not on q. The mean value of this quantity in the state considered is /= J/ Y»fo, *)/*•(&*) dfld*. (14.1) We introduce the function p(x', x) defined by 9 {x\ x) = j Y*(2, x')W(q, x) dq, (14.2) where the integration is extended only over the co-ordinates q ; this function is called the density matrix of the system. From the definition (14.2) it is evident that the function is "Hermitian": P *(x,x') = P (x',x). (14.3) The "diagonal elements" of the density matrix p(x,x) = j\¥(q t x)\*6q evidently determine the probability distribution for the co-ordinates of the system. f In order that W(q, x) should (at a given instant) fall into such a product, the measurement as a result of which this state was brought about must completely describe the system considered and the remainder of the closed system separately. In order that Wiq, x) should continue to have this form at subsequent instants, it is necessary in addition that these parts of the closed system should not interact (see §2). Neither of these conditions is now assumed. §14 The density matrix 39 Using the density matrix, the mean value / can be written in the form ?= j\j P (x',x)] x >= x dx. (14.4) Here / acts only on the variables x in the function p(x', x) ; after calculating the result of its action, we put x' = x. We see that, if we know the density matrix, we can calculate the mean value of any quantity characterising the system. It follows from this that, by means of p(x', x), we can also determine the probabilities of various values of the physical quantities in the system. Thus we reach the conclusion that the state of a system which does not have a wave function can be described by means of a density matrix. This does not contain the co-ordinates q which do not belong to the system concerned, though, of course, it depends essentially on the state of the closed system as a whole. The description by means of the density matrix is the most general form of quantum-mechanical description of the system. The description by means of the wave function, on the other hand, is a particular case of this, cor- responding to a density matrix of the form p{x' y x) = ¥*(*')¥(*). The following important difference exists between this particular case and the general one.f For a state having a wave function there is always a complete set of measuring processes such that they lead with certainty to definite results (mathematically, this means that T is an eigenfunction of some opera- tor). For states having only a density matrix, on the other hand, there is no complete set of measuring processes whose result can be uniquely predicted. Let us now suppose that the system is closed, or became so at some instant. Then we can derive an equation giving the change in the density matrix with time, similar to the wave equation for the Y function. The derivation can be simplified by noticing that the required linear differential equation for p(x', x, t) must be satisfied in the particular case where the system has a wave function, i.e. P (x',x,t)=W*(x',tyF(x,t). Differentiating with respect to time and using the wave equation (8.1), we have dp dY(x,t) W(x',t) ih~ = ihY*(x\ t) — K -LL+ifW(x, t) — dt dt dt = Y»(*\ *)#¥(*, *)-¥(*, *)#'*Y*(*', t), where A is the Hamiltonian of the system, acting on a function of *, and A' is the same operator acting on a function of x'. The functions ¥*(*', t) and Y(#, t) can obviously be placed behind the respective operators tt and U\ and we thus obtain the required equation : ih dp{x', x, t)JBt = (fi-fi'*)p(x',x, t). (14.5) f States having a wave function are sometimes called "pure" states, as distinct from "mixed" states, which are described by a density matrix. 40 Energy and Momentum §14 Let W n (x, t) be the wave functions of the stationary states of the system, i.e. the eigenfunctions of its Hamiltonian. We expand the density matrix in terms of these functions ; the expansion consists of a double series in the functions T w (#, t) and ^¥ n (x', t), which we write in the form 9 {x\ x, t) = SS a wn Y M *(*', t)YJx, t) = SS a mn ifj n *(x')tf; m (x)^f^ E n-E^ (i4 <6 ) tn n For the density matrix, this expansion plays a part analogous to that of the expansion (10.3) for wave functions. Instead of the set of coefficients a n , we have here the double set of coefficients a mn . These clearly have the pro- perty of being "Hermitian", like the density matrix itself: a nm * = a mn . (14.7) For the mean value of some quantity/ we have, substituting (14.6) in (14.4), / = SS a mn j Y n *(x, t)f<¥ m {x, t) dx, / = SZ a mn f n Jt) = SS a mn f nm &/ME n -E m )t } (14 . 8) or where f mn are the matrix elements of the quantity /. This expression is similar to formula (ll.l).f The quantities a mn must satisfy certain inequalities. The "diagonal elements" p(x, x) of the density matrix, which determine the probability distribution for the co-ordinates, must obviously be positive quantities. It therefore follows from the expression (14.6) (with x' = x) that the quadratic form constructed with the coefficients a nm (where the £ n are arbitrary complex quantities) must be positive. This places certain conditions, known from the theory of quadratic forms, on the quantities a nm . In particular, all the "diagonal" quantities must clearly be positive: a nn >0, (14.9) and any three quantities a nn , a mm and a mn must satisfy the inequality > Kn| 2 - (14.10) *nn'*mm To the "pure" case, where the density matrix reduces to a product of functions, there evidently corresponds a matrix a mn of the form We shall indicate a simple criterion which enables us to decide, from the f The description of a system by means of the quantities a mn was introduced independently by L. Landau and F. Bloch in 1927. §15 Momentum 41 form of the matrix %», whether we are concerned with a "pure" or a "mixed" state. In the pure case we have ( a )mn = ^ a mk a kn = S a k *a m a n *a k = a m a n * X\a k \ 2 k or (0 2 )«m = tf«m> (14.11) i.e. the density matrix is equal to its own square. §15. Momentum Let us consider a system of particles not in an external field. Since all positions in space of such a system as a whole are equivalent, we can say, in particular, that the Hamiltonian of the system does not vary when the system undergoes a parallel displacement over any distance. It is sufficient that this condition should be fulfilled for an arbitrary small displacement. An infinitely small parallel displacement over a distance Sr signifies a trans- formation under which the radius vectors r a of all the particles (a being the number of the particle) receive the same increment Sr : r a -> r -f- Sr. An arbitrary function tfjfa, r 2 , ... ) of the co-ordinates of the particles, under such a transformation, becomes the function ^(rj+Sr, r 2 +Sr, ... ) = ^x lt r 2 , ... )+Sr . S V a ^ = (l+8r.2Va¥(ri,r 2 ,...) a (V a denotes a "vector" whose components are the operators djdx a , d[dy a , djdz a ). The expression in parentheses, i.e. 1+Sr . 2 V«, a can be regarded as the operator of an infinitely small displacement, which converts the function «/r(r 1} r 2 , ... ) into the function ^fo+Sr, r 2 +Sr, ... ). The statement that some transformation does not change the Hamiltonian means that, if we make this transformation on the function i?«/r, the result is the same as if we make it only on the function if/ and then apply the operator i?. Mathematically, this can be written as follows. Let be the operator which effects the transformation in question. Then we have 0(fiip) = fi{Oift), whence 6ft-fi6 = o, i.e. the Hamiltonian must commute with the operator 0. In the case considered, the operator is the above-mentioned operator of an infinitely small displacement. Since the unit operator (the operator of multiplying by unity) commutes, of course, with any operator, and the 42 Energy and Momentum §15 constant factor Sr can be taken in front of 6, the condition OS—IiO = reduces here to (S V«)#-#(S V„) = 0. (15.1) a a As we know, the commutability of an operator (not containing the time explicitly) with i? means that the physical quantity corresponding to that operator is conserved. The quantity whose conservation for a closed system follows from the homogeneity of space is called momentum. Thus the relation (15.1) expresses the law of conservation of momentum in quantum mechanics ; the operator SV« must correspond, apart from a constant factor, to the total momentum of the system, and each term V a °f tne sum t° the momen- tum of an individual particle. The coefficient of proportionality between the operator p of the momentum of a particle and the operator V can be determined by means of the passage to the limit of classical mechanics. Putting p = c V and using the limiting expression (6.1) for the wave function, we have pT = (i/h) caF/MVS = c(i(hy¥S7S, i.e. in the classical approximation the effect of the operator p reduces to multiplication by (ijfijcS/S. The gradient S/S is, as we know from mech- anics, the momentum p of the particle; hence we must have (ilti)c= 1, i.e. c = — ih. Thus the operator of the momentum of a particle is p = —ih\J, or, in components, p x = —ikd/dx, p v = —ihd/dy, ft z = —ihdjdz. (15.2) It is easy to see that these operators are Hermitian, as they should be. For, with arbitrary functions ifs(x) and <f>(x) which vanish at infinity, we have <f>p x ip dx — —ih <f> — d# = ih ip — d# = *pp x *<f> d#, and this is the condition that the operator should be Hermitian. Since the result of differentiating functions with respect to two different variables is independent of the order of differentiation, it is clear that the operators of the three components of momentum commute with one another: W,-M- = o. AA-M. = °» AA-A?. = o- (is.3) This means that all three components of the momentum of a particle can simultaneously have definite values. Let us find the eigenf unctions and eigenvalues of the momentum operators. They are determined by the equations -ihdiMdx=p 3 xl>, -ihd/dy =p v f, -ihfy/dz = prf. (15.4) The solution of the first of these equations is §15 Momentum 43 where /is independent of x. This solution remains finite for all values of x, for any real value of p x . Thus the eigenvalues of the component p x of the momentum form a continuous spectrum extending from — oo to +00; the same is true, of course, of the components p y and p z . The three equations (15.4) have, in particular, common solutions, which correspond to states with definite values of all three momentum components forming the vector p. These solutions are of the form = C^/»>P- r , (15.5) where C is a constant. If all three components of the momentum are given simultaneously, we see that this completely determines the wave function of the particle. In other words, the quantities ^/ty, p z form one of the poss- ible complete sets of physical quantities. Let us determine the normalisation coefficient in (15.5). According to the rule (5.4) for normalising the eigenfunctions of a continuous spectrum, we must have J* <Mp* dV = S(p'- p) (15.6) (where dV = dxdydz), the integration being extended over all space; S(p'— p) is the three-dimensional delta function, defined similarly to the one-dimen- sional function (the delta function of one variable). f The integration can be immediately effected by means of the formulaj (1/2tt) J e l « x dx = 8(a). (15.7) We have f «£ p/ p * dV = C 2 I e<*/»XP'-P).r dV = C 2 (2^) 3 S(p'-p). Hence we see that we must have C^ilirKf = 1. Thus the normalised func- tion «/r is ^ p = (277-^)- 3 / 2 ^/»)P- r . (15.8) t The three-dimensional function S(r) can, in particular, be represented as a product of delta functions of the Cartesian components of the vector r: S(r) = 8(*)8(y)S(sr). I The conventional meaning of this formula is that the function on the left-hand side has the property of the delta function expressed by the equation J/(*)8(*)d*=/(0). This follows from the Fourier integral formula f(x') = (1/2tt) \\f{x)e* x ~ x '* dada, if we put *' = 0. Separating the real part, we can also write formula (15.7) in the form 00 (1/277-) J cosouc d;v = 8(a). (15.7a) 44 Energy and Momentum §15 The expansion of an arbitrary wave function ^(r) of a particle in terms of the eigenfunctions «/r p of its momentum operator is simply the expansion as a Fourier integral: 0(r) = j a(p)Mr) d 3 /> = (2^)- 3 /2 J a(p)e^l^- r d*p (15.9) (where d 3 /> = dp x dp y dp z ). The expansion coefficients a(p) are, according to formula (5.3), «(P) = J 0(r)0P*(r) dF = (2tt^)- 3 / 2 J 0(r)<r<W dF. (15.10) The function a(p) can be regarded (see §5) as the wave function of the particle in the "p representation"; \a(p)\ 2 d 3 p is the probability that the momentum has a value in the interval d 3 p. The formulae (15.9) and (15.10) give the relation between the wave functions in the two representations. Just as the operator p corresponds to the momentum, determining its eigenfunctions in the "r representation", we can introduce the idea of the operator f of the radius vector of the particle in the "p representation". It must be defined so that the mean value of the co-ordinates can be represented in the form r= J a*(p)*a(p) &p. (15.11) On the other hand, this mean value is determined from the wave function 0(r) by r = jif,*rjdV. Writing «/r(r) in the form (15.9) we have (integrating by parts)f r0(r) = (Zirh)- 3 ' 2 f ra(p)^/« p - r d^ = (2tt/*)- 3 / 2 f *M/«P- r [d a (p)/dp] &p. Using this expression and (15.10), we find f=j 0*r^ dV = (2ttH)- s / 2 j j 0*(r)#*[dtf(p)/dpy*"/A)p.r d 3 pdV = J iha*(p)[da(p)/dp] d*p. Comparing with (15.11), we see that the radius vector operator in the "p representation" is r = ihd/dp. (15.12) t The derivative with respect to the vector p is understood as the vector whose components are the derivatives with respect to p z , p y , p t . §15 Momentum 45 The momentum operator in this representation reduces simply to multipli- cation by p. PROBLEMS Problem 1. Express the operator T a of a parallel displacement over a finite distance a in terms of the momentum operator. Solution. By the definition of the operator T a we must have T*l>{r) = flr+a). Expanding the function ^(r+a) in a Taylor series, we have 0(r+a) = 0(r)+a . &£(r)/dr+ ... , or, introducing the operator p = —ih\J, «A(r+a) = £l+ia -P+~(^ -p) 2 + ••• ]tf(r). The expression in brackets is the operator This is the required operator of the finite displacement. Problem 2. Find the law of transformation of the wave function under a Galilean trans- formation. Solution. We shall carry out the transformation on a wave function of free motion of a particle (a plane wave). Since any function T can be expanded in terms of plane waves, this gives also the transformation law for an arbitrary wave function. The wave functions of free motion in frames of reference K and K' (where K' moves relative to K with velocity V) are T = constant xe ( */*> <!»•*-**>, Y' = constant x ««'»"»»'•*'-*'*>, with r = r'+\t, and the momenta and energies in the two systems are related byt p = p' + wV, E = E' + V.p'+^mV 2 . Substituting these expressions for r, p and £ in T, we obtain Y =Y' exp\-mV.(r' + iVt) 1 = T'expUmV.i(r+r')l. (1) In this form the formula involves no quantities characterising the free motion of the particle, and gives the required general law of transformation of the wave function of an arbitrary state of the particle. For a system of particles a sum over the particles appears in the ex- ponent in (1). t See Mechanics, 46 Energy and Momentum §16 §16. Uncertainty relations Let us derive the rules for commutation between momentum and co- ordinate operators. Since the result of successively differentiating with respect to one of the variables x, y, z and multiplying by another of them does not depend on the order of these operations, we have Pxy~ypx = 0, p x z-zp x = 0, (16.1) and similarly for P v , ft z . To derive the commutation rule for p x and x, we write (ftxX—xfixJtp = —ih d(xifi)ldx+ihx difi/dx = —ihifj. We see that the result of the action of the operator $xX—xp x reduces to multiplication by — ih; the same is true, of course, of the commutation of p y with y and p s with z. Thus we havef p x x—xp x = —ih, p v y—yp y = —ih, p z z—zp g = —ih. (16.2) All the relations (16.1) and (16.2) can be written in the form fiiX k —x k fii = —ihS ik (i,k = x,y,z). (16.3) Before going on to examine the physical significance of these relations and their consequences, we shall derive two formulae which will be useful later. Let /(r) be some function of the co-ordinates. Then p/(r)-/(r)P = -«»V/. (16.4) For (p/-/p)<A = -«[V(#)-/V0] - -«tyV/. A similar relation holds for the commutation of r with a "function" /(p) of the momentum operator: /(P)r-r/(p) = -ihdfldp. (16.5) It can be derived in the same way as (16.4) if we calculate in the p representa- tion, using the expression (15.12) for the co-ordinate operators. The relations (16.1) and (16.2) show that the co-ordinate of a particle along one of the axes can have a definite value at the same time as the components of the momentum along the other two axes ; the co-ordinate and momentum component along the same axis, however, cannot exist simultaneously. In particular, the particle cannot be at a definite point in space and at the same time have a definite momentum p. Let us suppose that the particle is in some finite region of space, whose dimensions along the three axes are (of the order of magnitude of) A#, Ay, Az. t These relations, discovered in matrix form by Heisenberg in 1925, formed the genesis of modern quantum mechanics. §16 Uncertainty relations 47 Also, let the mean value of the momentum of the particle be p . Mathe- matically, this means that the wave function has the form ijs = M(r)e< f /A > p «* r > where u(t) is a function which differs considerably from zero only in the region of space concerned. We expand the function ip in terms of the eigen- functions of the momentum operator (i.e. as a Fourier integral). The co- efficients a(p) in this expansion are determined by the integrals (15.10) of functions of the form u(r)e< i/n ^ p '~ p >- r . If this integral is to differ consider- ably from zero, the periods of the oscillatory factor g(*/»)(Po-p)- r must not be small in comparison with the dimensions Ax, Ay, Az of the region in which the function u(r) is different from zero. This means that a(p) will be con- siderably different from zero only for values of p such that (llh)(p 0x —p x )Ax <; 1, etc. Since |«(p)| 2 determines the probability of the various values of the momentum, the ranges of values oi p x , p y , p z m which a(p) differs from zero are just those in which the components of the momentum of the particle may be found, in the state considered. Denoting these ranges by Ap x , Ap v , Ap et we thus have ApJ^x ~ h, Ap y Ay ~ h, Ap z Az ~ h. (16.6) These relations, known as the uncertainty relations, were obtained by Heisenberg. We see that, the greater the accuracy with which the co-ordinate of the particle is known (i.e. the less Ax), the greater the uncertainty Ap x in the component of the momentum along the same axis, and vice versa. In parti- cular, if the particle is at some completely definite point in space (Ax = Ay = Az = 0), then Ap x = Ap y = Ap z = co. This means that all values of the momentum are equally probable. Conversely, if the particle has a completely definite momentum p, then all positions of it in space are equally probable (this is seen directly from the wave function (15.8), whose squared modulus is quite independent of the co-ordinates). As an example, let us consider a particle in a state described by the wave function ift = constant x (MUM**-*** I™ (16.7) (for simplicity, we consider a one-dimensional case, with the wave function depending on only one co-ordinate). The probabilities of the various values of the co-ordinates are |0| 2 = constant x e~ aX *l n , i.e. are distributed about the origin of co-ordinates (the mean value x = 0) according to a Gaussian law, with a standard deviation \Z[(AxY] = -\/(^/2a) (Ax denotes the difference x— x)-\. If the expansion coefficients a(p x ) of f As is well known, the Gaussian distribution for the probability w(x) of the values of some quantity x has the form w(x) = [2 7 r(Aj ( ;)2]-l/2 e -(A a; )V2(Ax) 2 48 Energy and Momentum §16 this function are calculated as a Fourier integral according to the formula a(p x ) = (2irA)-i/2 j ^( x )e-a/h)P x x dX) we obtain an expression of the form a(p x ) = constant x e-fcv-pJ'/M**. The distribution of probabilities of values of the momentum is \a\ 2 = constant X erl'i-'x)*/**, an( j consequently is also of Gaussian form, with a standard deviation (where &p x =p x —p )- The product of the standard deviations of co- ordinate and momentum is thus V[(A/0 2 (A*) 2 ]=p, (16.8) in agreement with the relation (16.6).f Finally, we shall derive another useful relation. Let / and g be two physical quantities whose operators obey the commutation rule M-gf=-ihc, (16.9) f It can be shown that this value of the product of the standard deviations is the least possible. To do this, we give the following formal derivation (H. Weyl). Let the state of the particle be de- scribed by the function *ji(x); for simplicity, we suppose the mean values of co-ordinate and momen- tum in this state to be zero. We consider the obvious inequality I dx dx >0, where a is an arbitrary real constant (the equality sign holds for a function of the form (16.7), and for no other). On calculating this integral, noticing that ! we obtain * 2 |0| 2 d* = (A*) 2 , Kx— f+xif,*— ) dx = f *— dx = - f M 2 dx = - 1, dx dx/ J dx J r d0* diL r d 2 <A 1 r „ 1 a 2 (A*) 2 -a+(l//* 2 )(A/g 2 > 0. If this quadratic (in a) trinomial is positive for all a, the condition 4(A*) 2 (1//* 2 )(A^) 2 >1 V[(A*) 2 (A/>x) 2 ]>P (16.8a) must be fulfilled. §16 Uncertainty relations 49 where 6 is the operator of some physical quantity c. On the right-hand side of the equation the factor h is introduced in accordance with the fact that in the classical limit (i.e. as h -» 0) all operators of physical quantities reduce to multiplication by these quantities and commute with one another. Thus, in the "quasi-classical" case, we can, to a first approximation, regard the right- hand side of equation (16.9) as being zero. In the next approximation, the operator c can be replaced by the operator of simple multiplication by the quantity c. We then have f£-£f= -fa- This equation is exactly analogous to the relation pxX— x Px = — ih, the only difference being that, instead of the constant h, we havef the quantity he. We can therefore conclude, by analogy with the relation AxAp x ~ h, that in the quasi-classical case there is an uncertainty relation A/<V ~ he (16.10) for the quantities / and g. In particular, if one of these quantities is the energy (/=i?) and the operator (§) of the other does not depend explicitly on the time, then by (9.2) c = g, and the uncertainty relation in the quasi-classical case is &EAg~hg. (16.11) f The classical quantity c is the Poisson bracket of the quantities / and g; see the footnote in §9. CHAPTER III SCHRODINGER'S equation §17. Schrodinger's equation Let us now turn to determining the form of the Hamiltonian — a problem of the greatest importance, since its solution determines the form of the wave equation. We shall begin by considering one free particle, i.e. a particle which is not in any external field. Because of the complete homogeneity of space for such a particle, its Hamiltonian cannot explicitly contain the co-ordinates, and must be expressible in terms of the momentum operator only. Moreover, for a free particle both the energy and the momentum are conserved, and hence both these quantities can exist simultaneously. Since the value of the momentum vector completely determines the state of the particle, the eigen- values of the energy E must be expressible in the form of functions of the value of the momentum in the same state. Here E is a function only of the absolute value of the momentum, and not of its direction ; this follows from the complete isotropy of space relative to the free particle, i.e. the equivalence of all directions in space. The actual form of the function E(p) is completely determined by the requirements of what is called Galileo's relativity principle, which must hold in non-relativistic quantum mechanics just as much as in classical (non-relativistic) mechanics. As is found in mechanics,! this re- quirement leads to a quadratic dependence of the energy on the momentum : E = p 2 j2m, where the constant m is called the mass of the particle. If the relation E = p 2 j2m holds for every eigenvalue of the energy and momentum, the same relation must hold for their operators also: # = (l/2«)tf.«+/,»+M (17.1) Substituting here from (15.2), we obtain the Hamiltonian of a freely moving particle in the form #=-(#/2m)A, (17.2) where A = d 2 / dx 2 + d 2 j dy 2 + d 2 jdz 2 is the Laplacian operator. If we have a system of non-interacting particles, its Hamiltonian is equal to the sum of the Hamiltonians of the separate particles : #=-P*S(l/m a )A (17.3) a (the suffix a is the number of the particle; A a is the Laplacian operator in f See Mechanics, §4. 50 §17 Schrddinger's equation 51 which the differentiation is with respect to the co-ordinates of the ath particle). The form of the Hamiltonian for a system of particles which interact with one another cannot be derived from the general principles of quantum mech- anics alone. It is found that it has in fact a form similar to that of Hamilton's function in classical mechanics: it is obtained by adding to the Hamiltonian of the non-interacting particles a certain function U(r ly r 2 , ... ) of their co- ordinates : #=-**■ S A^m a +Z/(r 1 ,r lf ...). (17.4) a The first term can be regarded as the operator of the kinetic energy and the second as that of the potential energy. The latter reduces to simple multipli- cation by the function U, and it follows from the passage to the limiting case of classical mechanics that this function must coincide with the one which gives the potential energy in classical mechanics. In particular, the Hamil- tonian for a single particle in an external field is ti =p 2 /2m+ U(x,y, z) = -(h*/2m) A + U(x,y, *), (17.5) where U(x, y, z) is the potential energy of the particle in the external field. The eigenvalues of the kinetic energy operator are positive; this follows at once from the fact that this operator is equal to the sum of the squares of the operators of the momentum components with positive coefficients. Hence the mean value of the kinetic energy in any state is also positive. Substituting the expressions (17.2) to (17.5) in the general equation (8.1), we obtain the wave equations for the corresponding systems. We shall write out here the wave equation for a particle in an external field: ih BV/dt = -(h*/2m)A x ¥+ U{x,y, zj¥. (17.6) The equation (10.2), which determines the stationary states, takes the form (h*/2m)A*fi +[E- U(x,y, z)]f = 0. (17.7) The equations (17.6) and (17.7) were obtained by Schrodinger in 1926 and are called Schrddinger's equations, with and without the time respectively. For a free particle, Schrodinger's equation (17.7) has the form (h*/2m)A<P+E+ = 0. (17.8) This equation has solutions finite in all space for any positive (or zero) value of the energy E. These solutions can be taken to be the common eigenfunc- tions (15.5) of the operators of the three momentum components. The com- plete wave functions of the stationary states will then have the form ¥ = constant x^/^W/fllP-r (g _ pZj2 m ). (17.9) Each such function describes a state in which the particle has a definite energy E and momentum p. This is a plane wave propagated in the direction 3 52 Schrodinger's Equation §17 of p and having an angular frequency E\h and wavelength 2irh\p (the latter is called the de Broglie wavelength of the particle).-)- The energy spectrum of a freely moving particle is thus found to be con- tinuous, extending from zero to + 00. Each of these eigenvalues (except E = 0) is degenerate, and the degeneracy is infinite. For there corresponds to every value of E, different from zero, an infinite number of eigenfunctions (17.9), differing in the direction of the vector p, which has a constant absolute magnitude. Let us enquire how the passage to the limit of classical mechanics occurs in Schrodinger's equation, considering for simplicity only a single particle in an external field. Substituting in Schrodinger's equation (17.6) the limit- ing expression (6.1) for the wave function, T = fle< i/a > s , we obtain, on per- forming the differentiation, dS da a ih . ih n h? a ih— +— (VS) 2 aAS \/S . \/a Aa+C/a = 0. dt Bt 2m 2m m 2m In this equation there are purely real and purely imaginary terms (we recall that S and a are real); equating each separately to zero, we obtain two equations dS 1 * a a — +-(\/S)*+U-— Aa = 0, dt 2m 2ma da a . 1 _ + _AS+— S/S . Vfl = 0. dt 2m m Neglecting the term containing h 2 in the first of these equations, we obtain -^+i-(V5) 2 +t/ = 0, (17.10) dt 2m that is, the familiar classical Hamilton- Jacobi equation for the action S of a particle, as it should be. We see, incidentally, that, as h -> 0, classical mech- anics is valid as far as quantities of the first (and not only the zero) order in h inclusive. The second equation obtained above, on multiplication by 2a, can be re- written in the form ™ +di J a *El)=0. (17.11) dt \ m J This equation has an obvious physical meaning: a 2 is the probability density for finding the particle at some point in space (|T*| 2 = a 2 ); VSj/i = p/p is the classical velocity v of the particle. Hence equation (17.11) is simply the equation of continuity, which shows that the probability density "moves" according to the laws of classical mechanics with the classical velocity v at every point. f The idea of a wave related to a particle was first introduced by L. db Broglie in 1924. §18 The fundamental properties of Schrodinger's equation 5 3 §18. The fundamental properties of Schrodinger's equation The conditions which must be satisfied by solutions of Schrodinger's equation are very general in character. First of all, the wave function must be single-valued and continuous in all space. The requirement of continuity is maintained even in cases where the field U(x, y, z) itself has a surface of discontinuity. At such a surface both the wave function and its derivatives must remain continuous. Concerning the continuity of the derivatives, how- ever, it must be added that this does not hold if there is some surface beyond which the potential energy U becomes infinite. A particle clearly cannot penetrate at all into a region of space where U = oo, i.e. we must have if/ = everywhere in this region. The continuity of if* means that ifi vanishes at the boundary of this region; the derivatives of i(j, however, in general are discontinuous in this case. If the field U(x, j, z) nowhere becomes infinite, then the wave function also must be finite in all space. The same condition must hold in cases where U becomes infinite at some point but does so only as l/r s with s < 2 (see also §35). Let U min be the least value of the function U(x, y, z). Since the Hamil- tonian of a particle is the sum of two terms, the operators of the kinetic energy (T ) and of the potential energy, the mean value E of the energy in any state is equal to the sum T+ V. But all the eigenvalues of the operator t (which is the Hamiltonian of a free particle) are positive; hence the mean value T > 0. Recalling also the obvious inequality V > £/ min , we find that & > ^min- Since this inequality holds for any state, it is clear that it is valid for all the eigenvalues of the energy: E n>U min . (18.1) Let us consider a particle moving in an external field which vanishes at infinity ; we define the function U(x, y y z), in the usual way, so that it vanishes at infinity. It is easy to see that the spectrum of negative eigenvalues of the energy will then be discrete, i.e. all states with E < in a field which vanishes at infinity are bound states. For, in the stationary states of a continuous spectrum, which correspond to infinite motion, the particle reaches infinity (see §10); however, at sufficiently large distances the field may be neglected, the motion of the particle may be regarded as free, and the energy of a freely moving particle can only be positive. The positive eigenvalues, on the other hand, form a continuous spectrum and correspond to an infinite motion; for E > 0, Schrodinger's equation in general has no solutions (in the field concerned) for which the integral / \if>\ 2 dV converges. f Attention must be drawn to the fact that, in quantum mechanics, a particle in a finite motion may be found in those regions of space where E < U; t However it must be mentioned that, for some particular mathematical forms of the function U(*, y, z) (which have no physical significance), a discrete set of values may be absent from the otherwise continuous spectrum. 54 Schrddinger's Equation §18 the probability |«/r| 2 of rinding the particle tends rapidly to zero as the distance into such a region increases, yet it differs from zero at all finite distances. Here there is a fundamental difference from classical mechanics, in which a particle cannot penetrate into a region where U > E. In classical mechanics the impossibility of penetrating into this region is related to the fact that, for E < U, the kinetic energy would be negative, that is, the velocity would be imaginary, which is meaningless. In quantum mechanics, the eigen- values of the kinetic energy are likewise positive; nevertheless, we do not reach a contradiction here, since, if by a process of measurement a particle is localised at some definite point of space, the state of the particle is changed, as a result of this process, in such a way that it ceases in general to have any definite kinetic energy. If U(x,y, z) > in all space (and U -» at infinity), then, by the inequality (18.1), we have E n > 0. Since, on the other hand, for E > the spectrum must be continuous, we conclude that, in this case, the discrete spectrum is absent altogether, i.e. only an infinite motion of the particle is possible. Let us suppose that, at some point (which we take as origin), U tends to — oo in the manner U « -ocr- s (a > 0). (18.2) We consider a wave function finite in some small region (of radius r ) about the origin, and equal to zero outside this region. The uncertainty in the values of the co-ordinates of a particle in such a wave packet is of the order of r ; hence the uncertainty in the value of the momentum is ~ hjr Q . The mean value of the kinetic energy in this state is of the order of h 2 jmr 2 , and the mean value of the potential energy is ~ — a/r *. Let us first suppose that s > 2. Then the sum h 2 lmr 2 — a/r s takes arbitrarily large negative values for sufficiently small r . If, however, the mean energy can take such values, this always means that the energy has negative eigenvalues which are arbitrarily large in absolute value. The mo- tion of the particle in a very small region of space near the origin corresponds to the energy levels with large \E\. The "normal" state corresponds to a particle at the origin itself, i.e. the particle "falls" to the point r = 0. If, however, s < 2, the energy cannot take arbitrarily large negative values. The discrete spectrum begins at some finite negative value. In this case the particle does not ' 'fall" to the centre. It should be mentioned that, in classical mechanics, the "fall" of a particle to the centre would be possible in principle in any attractive field (i.e. for any positive s). The case 5 = 2 will be specially considered in §35. Next, let us investigate how the nature of the energy spectrum depends on the behaviour of the field at large distances. We suppose that, as r -* oo, the potential energy, which is negative, tends to zero according to the power law (18.2) (r is now large in this formula), and consider a wave packet "filling" a spherical shell of large radius r and thickness Ar <^ r . Then the order §19 The current density 55 of magnitude of the kinetic energy is again £ 2 /m(Ar) 2 , and of the potential energy, — <x/r *. We increase r , at the same time increasing Ar, in such a way that Ar increases proportionally to r . If s < 2, then the sum h 2 jm(Ar) 2 — <x/r * becomes negative for sufficiently large r . Hence it follows that there are stationary states of negative energy, in which the particle may be found, with a fair probability, at large distances from the origin. This, however, means that there are levels of arbitrarily small negative energy (it must be recalled that the wave functions rapidly tend to zero in the region of space where U > E). Thus, in this case, the discrete spectrum contains an infinite number of levels, which become denser and denser towards the level E = 0. If the field diminishes as — \\r* at infinity, with s > 2, then there are not levels of arbitrarily small negative energy. The discrete spectrum terminates at a level with a non-zero absolute value, so that the total number of levels is finite. Schrodinger's equation (without the time) is real, as are the conditions imposed on its solution. Hence its solutions ifs can always be taken as real.f The eigenfunctions of non-degenerate values of the energy are automatically real, apart from the unimportant phase factor. For ifj* satisfies the same equation as tfi, and therefore must also be an eigenfunction for the same value of the energy; hence, if this value is not degenerate, i/t and «/»* must be essen- tially the same, i.e. they can differ only by a constant factor (of modulus unity). The wave functions corresponding to the same degenerate energy level need not be real, however, but by a suitable choice of linear combinations of them we can always obtain a set of real functions. The complete wave functions Y are determined by an equation in whose coefficients i appears. This equation, however, retains the same form if we replace t in it by — t and at the same time take the complex conjugate. % Hence we can always choose the functions Y in such a way that Y and Y* differ only by the sign of the time, a result which we know already from formulae (10.1) and (10.3). As is well known, the equations of classical mechanics are unchanged by time reversal, i.e. when the sign of the time is reversed. In quantum mechanics, the symmetry with respect to the two directions of time is expressed, as we see, in the invariance of the wave equation when the sign of t is changed and Y is simultaneously replaced by Y*. However, it must be recalled that this symmetry here relates only to the equation, and not to the concept of measurement itself, which plays a fundamental part in quantum mechanics (as we have explained in detail in §7). §19. The current density In classical mechanics, the velocity of a particle is equal to its momentum divided by its mass. We shall show that the same relation holds in quantum mechanics, as we should expect. t These assertions are not valid for systems in a magnetic field (see Chapter XV). | It is assumed that the potential energy U does not depend explicitly on the time: the system is either closed or in a constant (non-magnetic) field. 56 Schrodinger *s Equation §19 According to the general formula (9.2) for the differentiation of operators with respect to time, we have for the velocity operator v = f v = (Hh){8t-tfl). Using the expression (17.5) for fi and formula (16.5), we obtain v=p/m. (19.1) Similar relations will clearly hold between the eigenvalues of the velocity and momentum, and between their mean values in any state. The velocity, like the momentum of a particle, cannot have a definite value simultaneously with the co-ordinates. But the velocity multiplied by an infinitely short time interval dt gives the displacement of the particle in the time dt. Hence the fact that the velocity cannot exist at the same time as the co-ordinates means that, if the particle is at a definite point in space at some instant, it has no definite position at an infinitely close subsequent instant. We may notice a useful formula for the operator / of the derivative, with respect to time, of some quantity /(r) which is a function of the radius vector of the particle. Bearing in mind that/ commutes with U(r) t we find /= (i/k)(8f-ffi) = (il2mh)(&f-m. Using (16.4), we can write p 2 / = P • (/P-^ V/) = p/ • P-ihP . V/, /P 2 = (P/+*W/).P = p/.p+^V/.p. Substituting in the formula for/, we obtain the required expression: /= (l/2m)(£ . V/+V/-P). (19-2) Next, let us find the acceleration operator. We have v = (P)(#v - v#) = (i/mh)(fip - p#) = (ilmh)(Up -pU) (all the terms in i? except U(r) commute with p). Using formula (16.4), we find mfi = -\/U. (19.3) This operator equation is exactly the same in form as the equation of motion (Newton's equation) in classical mechanics. The integral / |T| 2 dV, taken over some finite volume V, is the probability of finding the particle in this volume. Let us calculate the derivative of this probability with respect to time. We have j /. f / gvp** d x ¥\ i r §19 The current density 57 Substituting here fi = 8* = -(#/2m) A + U(x,y, z) and using the identity T FA X F*- X F*AT = div (TV^-T* VY), we obtain v where i denotes the vector — JVpdF=- fdividF, i = (^/2m)(YV x F*-T*V 1 F). (19.4) The integral of div i can be transformed by Gauss's theorem into an integral over the closed surface S which boundsf the volume V: d It JV|»dF=- fi.df. (19.5) It is seen from this that the vector i may be called the probability current density vector. The integral of this vector over a surface is the probability that the particle will cross the surface during unit time. The vector i and the prob- ability density |T| 2 satisfy the equation apF|2/dH-divi = 0, (19.6) which is analogous to the classical equation of continuity. Introducing the momentum operator, we can write the vector i in the form i = (l/2/n)CFp*r*+*F*pF). (19 j) It is useful to show how the orthogonality of the wave functions of states with different energies follows immediately from Schrodinger's equation. Let tf> m and t// n be two such functions; they satisfy the equations -(h*/2m)A,l, m +Uf m = E m f mi -(*V2») A£»+ Ety n * = BJS. We multiply the first of these by n * and the second by i/, m and subtract corresponding terms ; this gives (£ m -2? n )«M«* = (£ 2 /2m)(«A ro A^ B *-&*A« = (k*l2m) div ty m V«A»*-^„* 7W. If we now integrate both sides of this equation over all space, the right-hand t The surface element df is denned as a vector equal in magnitude to the area d/ of the element and directed along the outward normal. 58 Schrodinger's Equation §20 side, on transformation by Gauss's theorem, reduces to zero, and we obtain whence, by the hypothesis E m # E n , there follows the required orthogonality relation §20. The variational principle Schrodinger's equation, in the general form Zfy = Etff, can be obtained from the variational principle 8 J" «/»*(#-£)«£ d ? = 0. (20.1) Since <p is complex, we can vary iff and $* independently. Varying «/r*, we have \ Stfj*(fi-E)ip dq = 0, whence, because Sip* is arbitrary, we obtain the required equation Biff = Eif/. The variation of if* gives nothing different. For, varying iff and using the fact that the operator i? is Hermitian, we have f i}j*(fi-E)8iJj dq = J 80(#*-£)«/<* dq = 0, from which we obtain the complex conjugate equation i?*«/r* = Eifs*. The variational principle (20.1) requires an unconditional extremum of the integral. It can be stated in a different form by regarding J? as a Lagran- gian multiplier in a problem with the conditional extremum requirement sJ^ifyd^O, (20.2) the condition being j ifjif,* dq = 1. (20.3) The least value of the integral in (20.2) (with the condition (20.3)) is the first eigenvalue of the energy, i.e. the energy E of the normal state. The func- tion ift which gives this minimum is accordingly the wave function «/»o of the normal state, f The wave functions ifs n (n > 0) of the other stationary states correspond only to an extremum, and not to a true minimum of the integral. In order to obtain, from the condition that the integral in (20.2) is a mini- mum, the wave function «/a and the energy E\ of the state next to the normal one, we must restrict our choice to those functions ifs which satisfy not only the f In the rest of this section we shall suppose the wave functions tf> to be real; they can always be so chosen (if there is no magnetic field). §20 The variational principle 59 normalisation condition (20.3) but also the condition of orthogonality with the wave function $ of the normal state: j i[n[/ dq = 0. In general, if the wave functions i(f Qy X , ... , if} n _ x of the first n states (arranged in order of in- creasing energy) are known, the wave function of the next state gives a mini- mum of the integral in (20.2) with the additional conditions J>d$ = l, J# ro d ? = (m = 0,1, 2,... ,fi-l). (20.4) We shall give here some general theorems which can be proved from the variational principle, f The wave function i/f of the normal state does not become zero (or, as we say, has no nodes) for any finite values of the co-ordinates. £ In other words, it has the same sign in all space. Hence, it follows that the wave functions i// n (n > 0) of the other stationary states, being orthogonal to tjt 0t must have nodes (if tfj n is also of constant sign, the integral j ift if/ n dq cannot vanish). Next, from the fact that i/i has no nodes, it follows that the normal energy level cannot be degenerate. For, suppose the contrary to be true, and let «Ao> to be two different eigenfunctions corresponding to the level E Q . Any linear combination Ci// -\-c't/j f will also be an eigenf unction ; but by choosing the appropriate constants c, c\ we can always make this function vanish at any given point in space, i.e. we can obtain an eigenfunction with nodes. If the motion takes place in a bounded region of space, we must have ifj = at the boundary of this region (see §18). To determine the energy levels, it is necessary to find, from the variational principle, the minimum value of the integral in (20.2) with this boundary condition. The theorem that the wave function of the normal state has no nodes means in this case that «/r does not vanish anywhere inside this region. We notice that, as the dimensions of the region containing the motion increase, all the energy levels E n decrease; this follows immediately from the fact that an extension of the region increases the range of functions which can make the integral a minimum, and consequently the least value of the integral can only diminish. The expression J* 0#0 dq = f [- S (£ 2 /2m a )«AA a 0+ Ut/fi] dq for the states of the discrete spectrum of a particle may be transformed into another expression which is more convenient in practice. In the first term of the ^integrand we write 4>Aa*l> = div a (0Va«A)-(Va«A) 2 . t The proof of theorems concerning the zeros of eigenfunctions (see also §21) is given by M. A. Lavrent'ev and L. A. Lyusternik, The Calculus of Variations (Kurs variatsionnogo ischisleniya), 2nd edition, chapter IX, Moscow 1950; R. Courant and D. Hilbert, Methods of Mathematical Physics, volume I, chapter VI, Interscience, New York 1953. % This theorem and its consequences are not in general valid for the wave functions of systems consisting of several identical particles (see footnote at the end of §63). 60 Schrodinger's Equation §21 The integral of div a (^Va^) over all space is transformed into an integral over an infinitely distant closed surface, and since the wave functions of the states of a discrete spectrum tend to zero sufficiently rapidly at infinity, this integral vanishes. Thus j 0f dq = j [S (h*l2ma)( Va<£) 2 + £ty 2 ] dq. (20.5) §21. General properties of motion in one dimension If the potential energy of a particle depends on only one co-ordinate (x), then the wave function can be sought as the product of a function of y and z and a function of x only. The former of these is determined by Schro- dinger's equation for free motion, and the second by the one-dimensional Schrodinger's equation -l+-[E-U(xM = Q. (21.1) dx z ft 2 Similar one-dimensional equations are evidently obtained for the problem of motion in a field whose potential energy is U(x, y, z) = Ui(x) + Uz{y) + U^{z), i.e. can be written as a sum of functions each of which depends on only one of the co-ordinates. In §§22-24 we shall discuss a number of actual examples of such "one-dimensional" motion. Here we shall obtain some general properties of the motion. We shall show first of all that, in a one-dimensional problem, none of the energy levels of a discrete spectrum is degenerate. To prove this, suppose the contrary to be true, and let ift x and ip 2 be two different eigenfunctions corresponding to the same value of the energy. Since both of these satisfy the same equation (21.1), we have or *lti"*l't~- l l t i" , l l i s= (th e prime denotes differentiation with respect to x). Integrating this relation, we find 0i'02-"~0i02' — constant. (21.2) Since ifs x = tfj 2 = at infinity, the constant must be zero, and so or ipx'tyi = 0«7^2- Integrating again, we obtain ^ = constant x ^ 2 » i.e. the two functions are essentially identical. The following theorem (called the oscillation theorem) may be stated for the wave functions iftjx) of a discrete spectrum. The function *jt n (x) correspond- ing to the (n+l)th eigenvalue E n (the eigenvalues being arranged in order of magnitude), vanishes n times (for finitef values of x). t If the particle can be found only on a limited segment of the x-axis, we must consider the zeros of *l> n (x) within that segment. §21 General properties of motion in one dimension 61 We shall suppose that the function U(x) tends to finite limiting values as x -> ± oo (though it need not be a monotonic function). We take the limiting value U(+ oo) as the zero of energy (i.e. we put U( + oo) = 0), and we denote U(—co) by U , supposing that U > 0. The discrete spectrum lies in the range of energy values for which the particle cannot move off to infinity; for this to be so, the energy must be less than both limiting values C/(±oo), i.e. it must be negative: E < 0, (21.3) and we must, of course, have in any case E > t/ min , i.e. the function U(x) must have at least one minimum with U mia < 0. Let us now consider the range of positive energy values less than U : 0<E<U . (21.4) In this range the spectrum will be continuous, and the motion of the particle in the corresponding stationary states will be infinite, the particle moving off towards x = + oo. It is easy to see that none of the eigenvalues of the energy in this part of the spectrum is degenerate either. To show this, it is sufficient to notice that the proof given above (for the discrete spectrum) still holds if the functions if/ v if> 2 are zero at only one infinity (in the present case they tend to zero as x -> — oo). For sufficiently large positive values of x, we can neglect U(x) in Schro- dinger's equation (21.1): 0" +(2»#W = 0. This equation has real solutions in the form of a stationary plane wave if* = a cos(fcc+8), (21.5) where a and 8 are constants, and the wave number k = pjh = ^(ImE)^. This formula determines the asymptotic form (for x ->+ oo) of the wave functions of the non-degenerate energy levels in the range (21.4) of the continuous spectrum. For large negative values of x, Schrodinger's equation is «£" -(2m/£ 2 )(£/ o -£)0 = 0. The solution which does not become infinite as x -> — oo is if> = be* x t where k = V[2m{U -E)]/h. (21.6) This is the asymptotic form of the wave function as x -> — oo. Thus the wave function decreases exponentially in the region where E < U. Finally, for E > U Q (21.7) the spectrum will be continuous, and the motion will be infinite in both directions. In this part of the spectrum all the levels are doubly degenerate. 62 Schrodinger's Equation §21 This follows from the fact that the corresponding wave functions are deter- mined by the second-order equation (21.1), and both of the two independent solutions of this equation satisfy the necessary conditions at infinity (whereas, for instance, in the previous case one of the solutions became infinite as x -> — oo, and therefore had to be rejected). The asymptotic form of the wave function as x -> + oo is if, = ai e ikx +a 2 e- ikx , (21.8) and similarly for x -> — oo. The term e ikx corresponds to a particle moving to the right, and e~ ikx corresponds to one moving to the left. Let us suppose that the function U(x) is even [U(—x) = U(x)]. Then Schrodinger's equation (21.1) is unchanged when the sign of the co- ordinate is reversed. It follows that, if tf/(x) is some solution of this equation, then iff(-x) is also a solution, and coincides with ip(x) apart from a constant factor: if*(—x) = ap(x). Changing the sign of x again, we obtain if/(x) = c 2 ifj(x), whence c = ±1. Thus, for a potential energy which is symmetrical (relative to x = 0), the wave functions of the stationary states must be either even [ip( — x) = ifi(x)] or odd [ifi — {x) = — ^(*)].f In particular, the wave function of the ground state is even, since it cannot have a node, while an odd function always vanishes for x — [^(0) = —$(0) = 0]. To normalise the wave functions of one- dimensional motion (in a continu- ous spectrum), there is a simple method of determining the normalisation coefficient directly from the asymptotic expression for the wave function for large values of \x\. Let us consider the wave function of a motion infinite in one direction, i.e. of a stationary state in the range (21.4) of the continuous spectrum. The normalisation integral diverges as x -> oo (as x -> — oo, the function decreases exponentially, so that the integral rapidly converges). Hence, to determine the normalisation constant, we can replace «/r by its asymptotic value (for large positive x), and perform the integration, taking as the lower limit any finite value of x, say zero ; this amounts to neglecting a finite quantity in comparison with an infinite one. We shall show that the wave function normalised by the delta function of p (the momentum of the particle at infinity) must have the asymptotic form (21.5) with a = ^(IJrrh), i.e. iftp « V( 2 M) cos(kx+8) = \/(l/27rA)[e« feB ^>+e-'**<-»>]. (21.9) Since we do not intend to verify the orthogonality of the functions corre- sponding to different^), on substituting the functions (21.9) in the normali- sation integral J $p*typ dx we shall suppose the momenta p to be arbitrarily close; we can therefore put 8 = 8' (in general 8 is a function of p). Next, we t In this discussion it is assumed that the stationary state is not degenerate, i.e. the motion is not infinite in both directions. Otherwise, when the sign of x is changed, two wave functions belonging to the energy level concerned may be transformed into each other. In this case, however, although the wave functions of the stationary states need not be even or odd, they can always be made so (by choosing appropriate linear combinations ofrthe original functions). §22 The potential well 63 retain in the integrand only those terms which diverge for p = p'\ in other words, we omit terms containing the factor e±*(fc+fc')£. Thus we obtain f i/j p *iP P ' dx = {Ijlrrh) ( j e«*'-*>* dx+ j «-«*'-*>* dA or f $f$ v > dx = (1/2tt£) f «<*'-*)* d*. This integral, however, is identical with the normalisation integral for the wave function of free motion tfj p = (27rA)-i/2e**« (21.10) which is normalised by the delta function of momentum (cf. (15.8)). The change to normalisation by the delta function of energy is effected, in accordance with (5.14), by multiplying i/j p by s/{dpjdE) = lj-y/v, where v is the velocity. Thus for free motion we have 4s E = {2irhv)-Vty* x . (21.11) The probability current density in this wave is ©|0*|2 = 1/2tt*. (21.12) Dividing the function (21.9) by \/v and using equation (21.12), we can formulate the following rule for the normalisation of the wave function for a motion infinite in one direction by the delta function of energy: having represented the asymptotic expression for the wave function in the form of a sum of two plane waves travelling in opposite directions, we must choose the normalisation coefficient in such a way that the probability current density in the wave travelling towards (or away from) the origin is I/IttH. Similarly, we can obtain an analogous rule for normalising the wave func- tions of a motion infinite in both directions. The wave function will be normalised by the delta function of energy if the sum of the probability cur- rents in the waves travelling towards the origin from x = + oo and x = — oo is H2ttH. §22. The potential well As a simple example of one-dimensional motion, let us consider motion in a square potential well, i.e. in a field where U(x) has the form shown in Fig. 1 : U(x) = for < x < a, U(x) = U f or x < and x > a. It is evident a priori that for E < U the spectrum will be discrete, while for E > U we have a continuous spectrum of doubly degenerate levels. In the region < x < a we have SchrSdinger's equation ^'+(2mlh z )Ei/ t = (22.1) (the prime denotes differentiation with respect to x) y while in the region 64 Schrddinger's Equation §22 uix) a Fig. 1 outside the well f +(2»/#)(£- UM = 0. (22.2) For # = and a; = a the solutions of these equations must be continuous together with their derivatives, while for x = ± oo the solution of equation (22.2) must remain finite (for the discrete spectrum when E < U , it must vanish). For E < U , the solution of equation (22.2) which vanishes at infinity is tfi = constant xe***, where k = V[(2mJh 2 )(U -E)]; (22.3) the signs — and + in the exponent refer to the regions x > a and * < respectively. The probability \ift\ 2 of finding the particle decreases exponen- tially in the region where E < U(x). Instead of the continuity of t// and tfj' at the edge of the potential well, it is convenient to require the continuity of iff and of its logarithmic derivative ip'/ift. Taking account of (22.3), we obtain the boundary condition in the form f/0 = =F*. (22.4) We shall not pause here to determine the energy levels in a well of arbitrary depth U (see Problem 2), and shall analyse fully only the limiting case of infinitely high walls (U -> oo). For U = oo, the motion takes place only between the points x = and x = a and, as was pointed out in §18, the boundary condition at these points must be # = 0. (22.5) (It is easy to see that this condition is also obtained from the general condition (22.4). For, when U -> oo, we have also k -> oo and hence tf/'ji// -+ oo; since ip' cannot become infinite, it follows that ift = 0.) We seek a solution of equation (22.1) inside the well in the form ift = c sin(foc+8), where k = </(2mEI&). (22.6) The condition ^ = for x = gives S = 0, and then the same condition for §22 The potential well 65 x = a gives sin ka = 0, whence ka — rnr, n being a positive integer,! or E n = (7T 2 ^ 2 /2ma 2 )» 2 , n = 1,2,3,.... (22.7) This determines the energy levels of a particle in a potential well. The normalised wave functions of the stationary states are *l>n = vWO sin(7rnxla). (22.8) From these results we can immediately write down the energy levels for a particle in a rectangular "potential box", i.e. for three-dimensional motion in a field whose potential energy U = QforO < x < a y <y <b,0 < z <c and U = oo outside this region. In fact, these levels are given by the sums ir 2 h 2 /n, 2 Wo 2 « 3 2 \ *w, = T-( ^-+1F + -T- ) (*» W2 ' n * = 1,2,3,...), (22.9) 128 2m \ a 2 6 2 c 2 / and the corresponding wave functions by the products / 8 Trn x 7rw 2 rm 3 0«i",«« = H~ sin — ^^sin— ^ysin — z. (22.10) 1 s ' V abc a b c It may be noted that the energy Eq of the ground state is, by (22.7) or (22.9), of the order of h 2 fml 2 , where / is the linear dimension of the region in which the particle moves. This result is in accordance with the uncertainty relation ; when the uncertainty in the co-ordinate is ~ /, the uncertainty in the momentum, and therefore the order of magnitude of the momentum itself, is ~ hjl. The corresponding energy is ~ (hjl) 2 lm. PROBLEMS Problem 1. Determine the probability distribution for various values of the momentum for the normal state of a particle in an infinitely deep square potential well. Solution. The coefficients a(p) in the expansion of the function tp t (22.8) in terms of the eigenfunctions (21.10) of the momentum are a(p) = tf/j*^ dx = sinf -* )«-&/*>»>* dx. J -\/{iTah) J \a / Calculating the integral and squaring its modulus, we obtain the required probability distri- bution : 47rh z a pa |a(/>)| 2 = cos 2 —. ' ^ ' (p 2 a 2 -7r 2 ^ 2 ) 2 2H Problem 2. Determine the energy levels for the potential well shown in Fig. 2. Solution. The spectrum of energy values E < U u which we shall consider, is discrete. In the region * < the wave function is = aefi't where k x = ^/[{ImJh^U^-E)}, f For « = 0we should have = identically. 66 Schrodinger's Equation u(x) §22 Us <A Fig. 2 while in the region x > a if, = c 2 e-"* x , where /c 2 = V[(2m/h 2 )(U 2 -E)]. Inside the well (0 < x < a) we look for >p in the form iff = c sin(foc+8), where k=^(2mE/h 2 ). The condition of the continuity of ifi'/*fi at the edges of the well gives the equations k cot 3 = k x = V[(2™lh 2 )U 1 -k 2 ], k cot(ka+S) = -k 2 = -^/[(2m/h 2 )U 2 -k 2 ] y or sin 8 = kh/ViZmUJ, sin(&*+8) = -kh/V(2mU 2 ). Eliminating S, we obtain the transcendental equation ka = rnr- sirr^kh/ V(2m C/j)] - sivr 1 [khl^{2m U 2 )] (1) (where n = 1, 2, 3, ... , and the values of the inverse sine are taken between and Jtt), whose roots determine the energy levels E = k 2 h z /2m. For each n there is in general one root; the values of n number the levels in order of increasing energy. Since the argument of the inverse sine cannot exceed unity, it is clear that the values of k can lie only in the range from to V(2mU 1 /h 2 ). The left-hand side of equation (1) increases monotonically with k, and the right-hand side decreases monotonically. Hence it is neces- sary, for a root of equation (1) to exist, that for k = V(2mU 1 /h i ) the right-hand side should be less than the left-hand side. In particular, the inequality ay/ilmU^fh >frr-sm-i</(UJUJ, (2) which is obtained for « = 1, is the condition that at least one energy level exists in the well. We see that for given and unequal U u U 2 there are always widths a of the well which are so small that there is no discrete energy level. For t/j = U 2 , the condition (2) is evidently always satisfied. For Ui = U 2 = U (a symmetrical well), equation (1) reduces to sm- 1 [hk/V(2mU )] = \{mr-hd). (3) Introducing the variable £ = \ka, we obtain for odd n the equation COS £ = ±yi, where y = {t\\a)^{2\mVJ) y (4) and those roots of this equation must be taken for which tan g > 0. For even n we obtain the equation sin $ = ±yg t (5) and we must take those roots for which tan £ < 0. The roots of these two equations deter- mine the energy levels E = 2£ 2 h 2 /ma 2 . The number of levels is finite when y ^ 0. §23 The linear oscillator 67 In particular, for a shallow well in which U < h 2 /ma 2 , we have y > 1 and equation (5) has no root. Equation (4) has one root (with the upper sign on the right-hand side), s = 1/y— l/2y 3 . Thus the well contains only one energy level, E £ U -(ma*l2HV)Uo*, which is near the top of the well. Problem 3. Determine the pressure exerted on the walls of a rectangular "potential box" by a particle inside it. Solution. The force on the the wall perpendicular to the #-axis is the mean value of the derivative -dH/da of the Hamilton's function of the particle with respect to the length of the box in the direction of the *-axis. The pressure is obtained by dividing this force by the area be of the wall. According to the formula derived in §11, Problem, the required mean value is found by differentiating the eigenvalue (22.9) of the energy. The result is pU) _ -nZffin^jtnaZbc. §23. The linear oscillator Let us consider a particle executing small oscillations in one dimension (what is called a linear oscillator). The potential energy of such a particle is well known to be %mco 2 x 2 , where co is, in classical mechanics, the character- istic (angular) frequency of the oscillations. Accordingly, the Hamiltonian of the oscillator is & =iP 2 /m+ \mco 2 x 2 . (23.1) Since the potential energy becomes infinite for x = ±00, while its least value (at x = 0) is zero, it is clear from general principles that the energy spectrum of the oscillator is discrete and the energy values are positive. Let us determine the energy levels of the oscillator, using the matrix methodf. We shall start from the "equations of motion" in the form (19.3) ; in this case they give x+co 2 x = 0. (23.2) In matrix form, this equation reads (x)mn+<*>*x mn = 0. For the matrix elements of the acceleration we have, according to (11.8), (*)«m = lco mn(x)mn = — ^mn^mn' Hence we obtain (w WB 2 -w 2 )* mM = 0. Hence it is evident that all the matrix elements x mn vanish except those for which to mn = at or co mn = - co. We number all the stationary states so that the frequencies ± co correspond to transitions n ->w+ 1, i.e. co n ,r^i = ± co. Then the only non-zero matrix elements are x n , n±1 . t This was done by Heisenberg in 1925, before Schrodinger's discovery of the wave equation. 68 Schrodinger's Equation §23 We shall suppose that the wave functions ift n are taken real. Since x is a real quantity, all the matrix elements x mn are real. The Hermitian condition (11.10) now shows that the matrix x mn is symmetrical: To calculate the matrix elements of the co-ordinate which are different from zero, we use the commutation rule &x— x& = —thjm, written in the matrix form (xx) mn -(xx) mn = -(ihlm)8 mn . By the matrix multiplication rule (11.12) we hence have for m = n i 2 ((*inlXnlXln—Xnlo»lnXm) = 2* 2 a>nlXnl 2 = —th/m. I • In this sum, only the terms with / = n ± 1 are different from zero, so that we have (Xn + l,nY-{n,Xn-lY = h/lmw. (23.3) From this equation we deduce that the quantities (#n+i,n)* form an arith- metic progression, which is unbounded above, but is certainly bounded below, since it can contain only positive terms. Since we have as yet fixed only the relative positions of the numbers n of the states, but not their abso- lute values, we can arbitrarily choose the value of n corresponding to the first (normal) state of the oscillator, and put this value equal to zero. Accordingly x _! must be regarded as being zero identically, and the application of equa- tions (23.3) with n = 0, 1, ... successively leads to the result (*n,n-i) 2 = nh\2moi. Thus we finally obtain the following expression for the matrix elements of the co-ordinate which are different from zero:f *«,«-! = *-l« = vW2»c). ( 23 - 4 ) The matrix of the operator tt is diagonal, and the matrix elements H nn are the required eigenvalues E n of the energy of the oscillator. To calculate them, we write = \m[ S ioi nl x nl iw ln x ln +oi 2 S x n ix ln ] =|wS(a> 2 +co n? 2 )^ n 2 . In the sum over /, only the terms with /= n±\ are different from zero; t We choose the indeterminate phases o„ (see the third footnote to §11) so as to obtain the plus sign in front of the radical in all the matrix elements (23.4). Such a choice is always possible for a matrix in which only those elements are different from zero which correspond to transitions between states with adjacent numbers. §23 The linear oscillator 69 substituting (23.4), we obtain E n = (n+$)ha>, n = 0,1,2 (23.5) Thus the energy levels of the oscillator lie at equal intervals of hot from one another. The energy of the normal state (« = 0) is \ha\ we call atten- tion to the fact that it is not zero. The result (23.5) can also be obtained by solving Schrodinger's equation. For an oscillator, this has the form dV 2w ~^-^<E-lmo^x^ = 0. (23.6) Here it is convenient to introduce, instead of the co-ordinate x, the dimension- less variable g by the relation £ = ^{moijh)x. (23.7) Then we have the equation r+[(2Elko)-£ 2 W = 0; (23.8) here the prime denotes differentiation with respect to £. For large £, we can neglect lEjhco in comparison with £ 2 ; the equation «/," = ^tf, has the asymptotic integrals = c ±«' (for differentiation of this function gives ifi" = ^j, on neglecting terms of order less than that of the term retained). Since the wave function fi must remain finite as g ->±oo, the index must be taken with the minus sign. It is therefore natural to make in equation (23.8) the substitution ^ = «" f8/2 *(£)• (23.9) For the function x{£) we obtain the equation (with the notation {lEjhoi) — 1 = 2n; since we already know that E > 0, we have n > — £) x"-2&'+2»x = 0, (23.10) where the function x must be finite for all finite £, and for $ -»± oo must not tend to infinity more rapidly than every finite power of $ (in order that the function tf> should tend to zero). Such solutions of equation (23.10) exist only for positive integral (and zero) values of n (see §a of the Mathematical Appendices); this gives the eigenvalues (23.5) for the energy, which we know already. The solutions of equation (23.10) corresponding to various integral values of n are x = con- stant xH n (£), where H n (g) are what are called Hermite polynomials; these are polynomials of the nth degree in £, defined by the formula H n (€) = (-l)V>(<r**)/d£ n . (23.11) 70 Schrodinger's Equation §23 Determining the constants so that the functions ijs n satisfy the normalisation condition oo j ifr n 2 (x) dx = 1, —GO we obtain (see (a. 7)) -r) on/, // e-m^y^H n (xV[mcolh]). (23.12) Thus the wave function of the normal state is O («) = (mwlTrh) 1 ^-™"**/™. (23.13) It has no zeros for finite x, which is as it should be. By calculating the integrals J ift n *p m € d£, we can determine the matrix ele- ments of the co-ordinate; this calculation leads, of course, to the same values (23.4). Finally, we shall show how the wave functions ip n may be calculated by the matrix method. We notice that, in the matrices of the operators £±icox, the only elements different from zero are (x— tW)„_ 1>n = —{x+io}x) nn _ x = —{^(liohn/m). (23.14) Using the general formula- (11.11), and taking into account the fact that ift-i = 0, we conclude that (£—ia}x)tfi = 0. After substituting the expression £ = —{(hjmjdldx, we obtain the equation difijdx = —(mcolh)xif/ Qt whose normalised solution is (23.13). And, since (£+ia)x)if/ n _ 1 = (x+iu)x) n3t ^_ x tfj n = i^(2coknlm)ift n , we obtain the recurrence formula ty n = ■y/(ml2o)hn)[—(hlm) d/dx+coxtyn^ i / a \ i d V(2«)\ dg J V(2«) d£ when this is applied n times to the function (23.13), we obtain the expression (23.12) for the normalised functions ip n - PROBLEMS Problem 1. Determine the probability distribution of the various values of the momentum for an oscillator. Solution. Instead of expanding the wave function of the stationary state in terms of the eigenfunctions of momentum, it is simpler in the case of the oscillator to start directly from §23 The linear oscillator 7 1 Schrodinger's equation in the "p representation". Substituting in (23.1) the co-ordinate operator £ = ihdjdp (15.12), we obtain the Hamiltonian in the p representation, GER'i d 2 a(/>) 2 dp 2 rnuPW' The corresponding Schrodinger's equation Ba(p) — Ea(p) for the wave function a(p) in the p representation is ( E -ty p)= °- This equation is of exactly the same form as (23.6); hence its solutions can be written down at once by analogy with (23.12) (replacing xVitno/h) in this formula by p/V(mtoft)). Thus we find the required probability distribution to be 2 n n\y{Trma>n) Problem 2. Determine the lower limit of the possible values of the energy of an oscillator, using the uncertainty relation (16.8a). Solution. We have for the mean value of the energy of the oscillator or, using the relation (16.8a), E > (Ap) 2 /2m+ma) 2 h 2 /8(Ap)\ On determining the minimum value of this expression (regarded as a function of (Ap) 2 ), we find the lower limit of the mean values of the energy, and therefore that of all possible values : E>ihco. Problem 3. Determine the energy levels for a particle moving in a field of potential enerev (Fig. 3) U(x) = A(e- %ax — 2e~ ax ) (P. M. Morse). ' Solution. The spectrum of positive eigenvalues of the energy is continuous (and the levels are not degenerate), while the spectrum of negative eigenvalues is discrete. Schrodinger's equation reads We introduce a new variable 2V(2mA) £ = e -ax Cf.k 72 Schrodinger's Equation §23 (taking values from to oo) and the notation (we consider the discrete spectrum, so that E<0) s = ^(—2mE)laLh, n = V(2mA)/*h-(s+$). (1) Schrodinger's equation then takes the form r v ,+ (-< n+s+$ s 2 ■> 0. As £ -> oo, the function $ behaves asymptotically as e ± if, while as £ -v it is proportional to £±*. From considerations of finiteness we must choose the solution which behaves as e~it as £ -> oo and as £ s as £ -» 0. We make the substitution = e -t'H s z»($) and obtain for w the equation £w"+(2*+ 1-|)«>'+««; = 0, (2) which has to be solved with the conditions that w is finite as £ -> 0, while as £ -> oo, w tends to infinity not more rapidly than every finite power of £. Equation (2) is the equation for a confluent hypergeometric function (see §d of the Mathematical Appendices) : to = F(-n,2s+l y g). A solution satisfying the required conditions is obtained for non-negative integral n (when the function F reduces to a polynomial). According to the definitions (1), we thus obtain for the energy levels the values -E n = A^\ a.h .(»+*) ]'■ V(2mA) where n takes positive integral values from zero to the greatest value for which \/(2mA)[a.h > n-\- J (so that the parameter s is positive in accordance with its definition). Thus the discrete spectrum contains only a limited number of levels. If V(2mA)JctH < %, there is no discrete spectrum at all. Problem 4. The same as Problem 3, but with U = — L7 /cosh 2 ax (Fig. 4). Solution. The spectrum of positive eigenvalues of the energy is continuous, while that of negative values is discrete ; we shall consider the latter. Schrodinger's equation is dV 2m/ _L + _(£+- cbc 2 h 2 \ cosb^aa:. Up \ sh 2 aa: / 0. obtaining §24 Motion in a homogeneous field 73 We put £ = tanh auc and use the notation e = V(-2w£)//wc, 2mC/ /a 2 ^2 = *(s+l), 5[ ( '- p ?]*[* t "-ii>- 1 This is the equation of the associated Legendre polynomials; it can be brought to hyper- geometric form by making the substitution «/- = (1— £ 2 ) e/2 w(£) and temporarily changing the variable to « = J(l— £): M(l-ttK' + (e+l)(l-2z/)«;'-( e -s)(e + s+i)a> = 0. The solution finite for £ = 1 (i.e. for * = oo) is If tfi remains finite for £ = — 1 (i.e. for x = — oo), we must have e—s = —n, where n = 0, 1, 2, ...; then F is a polynomial of degree n, which is finite for f = —1. Thus the energy levels are determined by s — € = «, or /* 2 <x 2 r .... //. 8mU n WC) 8m There is a finite number of levels, determined by the condition e > 0, i.e. « < $. §24. Motion in a homogeneous field Let us consider the motion of a particle in a homogeneous external field. We take the direction of the field as the axis of x; let F be the force acting on the particle in this field. In an electric field of intensity E, this force is F = eE, where e is the charge on the particle. The potential energy of the particle in the homogeneous field is of the form U = —Fx+ constant; choosing the constant so that U = for x = 0, we have U = —Fx. Schrodinger's equation for this problem is d 2 if,ldx 2 +(2m/h 2 )(E+Fx)if, = 0. (24.1) Since U tends to + oo as x -> — oo, and vice versa, it is clear that the energy levels form a continuous spectrum occupying the whole range of energy values £from -co to +oo. None of these eigenvalues is degenerate, and they correspond to motion which is finite towards x = — oo and infinite to- wards X = +00. Instead of the co-ordinate x, we introduce the dimensionless variable f = (x+ElF)(2mF/&)V*. (24.2) Equation (24.1) then takes the form *"+# = 0. (24.3) 74 Schrodinger's Equation §24 This equation does not contain the energy parameter. Hence, if we obtain a solution of it which satisfies the necessary conditions of finiteness, we at once have the eigenfunction for arbitrary values of the energy. The solution of equation (24.3) which is finite for all x has the form (see §b of the Mathematical Appendices) m = A®(-£) t (24.4) where <D(£) = cos(£tt 3 +«!) du \/tt J v is called the Airy function, while A is a normalisation factor which we shall determine below. As | -> — oo, the function ifj(g) tends exponentially to zero. The asymp- totic expression which determines «/r(|) for large negative values of £ is (see (b.4)) ' A(a ~2J^ eXp[ " fI ^ |3/2] ' (24 " 5) For large positive values of £, the asymptotic expression for ?/r(|) is (see (b.5))t </,(£) = AiJT^> sin(f^/ 2 +|7r). (24.6) Using the general rule (5.4) for the normalisation of eigenfunctions of a continuous spectrum, let us reduce the function (24.4) to the form normalised by the delta function of energy, for which jW(f ) d* = 8(E'-E). (24.7) In §21 we gave a simple method of determining the normalisation coefficient by means of the asymptotic expression for the wave functions. Following this method, we represent the function (24.6) as the sum of two travelling waves : 0(0 a^^- 1 /4exp(f[f^/ 2 -|7r])+|^- 1 / 4 exp(-t[f^/ 2 -i7r]). The probability current density v |^| 2 , calculated from each of these two terms, must be \\2rnh: «/[2(E+Fx)ltn](AI2pi*)* = A\2hFyi*l4m*l* = 1/2^, whence we find (2m) 1 /» A = — — - . (24.8) ^1/2^1/6^2/3 f It may be noted, by way of anticipation, that the asymptotic expressions (24.5) and (24.6) cor- respond to the quasi-classical expressions (47.1) and (47.4a) for the wave function. §25 The transmission coefficient 75 PROBLEM Determine the wave functions in thep representation for a particle in a homogeneous field. Solution. The Hamiltonian operator in the p representation is fl = p*/2m-ihF d/dp, so that Schrodinger's equation for the wave function a(p) has the form da fp % \ -ihF—+[- — E)a = Q. dp \2m J Solving this equation, we find the required functions a E (p) = (277-^F)-i/ 2 ^'/ft^^P-P 3 /6m). These functions are normalised by the condition ja E *(p)a E {p)dp = S(E'-E). §25. The transmission coefficient Let us consider the motion of particles in a field of the type shown in Fig. 5 : U(x) increases monotonically from one constant limit ( U = as x -> — oo) to another (U = U Q as x -> +oo). According to classical mech- anics, a particle of energy E < U moving in such a field from left to right, on reaching such a "potential wall", is "reflected" from it, and begins to move in the opposite direction ; if, however, E > U , the particle continues to move in its original direction, though with diminished velocity. In quantum mechanics, a new phenomenon appears : even for E > U Q , the particle may be "reflected" from the potential wall. The probability of reflection must in principle be calculated as follows. Fig. 5 Let the particle be moving from left to right. For large positive values of x, the wave function must describe a particle which has passed "above the wall" and is moving in the positive direction of x, i.e. it must have the asymp- totic form for * -> oo, « Ae ik * x , where k 2 = (l/h)y/[2m(E— U Q )] (25.1) and A is a constant. To find the solution of Schrodinger's equation which satisfies this boundary condition, we calculate the asymptotic expression for 76 Schrodinger's Equation §25 x -> — oo ; it is a linear combination of the two solutions of the equation of free motion, i.e. it has the form for x -» - oo, tjt « e ile * x +Be- ik * x t where ^ = ^/(2mE)/h. (25.2) The first term corresponds to a particle incident on the "wall" (we suppose if> normalised so that the coefficient of this term is unity) ; the second term represents a particle reflected from the "wall". The probability current density in the incident wave is k ly in the reflected wave k^B] 2 , and in the transmitted wave £ 2 I^I 2 - We define the transmission coefficient D of the par- ticle as the ratio of the probability current density in the transmitted wave to that in the incident wave : D = (*A)M». (25.3) Similarly we can define the reflection coefficient R as the ratio of the density in the reflected wave to that in the incident wave. Evidently R = 1— D: R = \B\ 2 = l-(k 2 lk 1 )\A\ 2 (25.4) (this relation between A and B is automatically satisfied). If the particle moves from left to right with energy E < U , then k 2 is purely imaginary, and the wave function decreases exponentially as # -» + oo. The reflected current is equal to the incident one, i.e. we have "total reflec- tion" of the particle from the potential wall. We emphasise, however, that in this case the probability of finding the particle in the region where E < U is still different from zero, though it diminishes rapidly as x increases. In the general case of an arbitrary stationary state (with energy E > C/o), the asymptotic form of the wave function is given, both for x ->— oo and for x -> + oo, by a sum of waves propagated in each direction: di = A\e iklX +B\er ilc ^ x for x -> — oo, (25.5) *fi = A2,e ik * x +B2fir ik * x for x -> +oo. Since these expressions are asymptotic forms of the same solution of a linear differential equation, there must be a linear relation between the coefficients Ai, B± and A2, B%. Let A% = a.Ai+f$B\, where a, j8 are constants (in general complex) which depend on the specific form of the field U(x). The corres- ponding relation for B2 can then be written down from the fact that Schro- dinger's equation is real. This shows that, if ip is a solution of a given Schrodinger's equation, the complex conjugate function iff* is also a solution. The asymptotic forms «/»* = A 1 *e~ ik ^ x + B±*e iklX for *-^— 00, tjj# = A 2 * e -ik2x + B 2 *e ik * x for *->+oo differ from (25.5) only in the nomenclature of the constant coefficients; we therefore have B 2 * = aBi*+fiAi* or B 2 = ol*Bi+P*Ai. Thus the coefficients §25 The transmission coefficient 77 in (25.5) are related by equations of the form A 2 = o^i+jSBi, j? 2 = jS*^i+a*#i. (25.6) The condition of constant probability current along the jc-axis leads to the relation k^A^-^) = k*{\A 2 \*-\B 2 \% Expressing A 2 , B 2 in terms of Ai, B\ by (25.6), we find |a| 2 -|j3| 2 = hlk 2 . (25.7) Using the relation (25.6), we can show, in particular, that the reflection coefficients are equal (for a given energy E > Uq) for particles moving in the positive and negative directions of the x-axis ; the former case corresponds to putting B 2 = in (25.5), and the latter case to A\ = 0. The corresponding reflection coefficients are Ri = |#i/^i| 2 = |£*/a*|2, R 2 = \A 2 jB 2 \* = |)3/a*|2, whence it is clear that R± = R 2 . PROBLEMS Problem 1. Determine the reflection coefficient of a particle from a rectangular potential wall (Fig. 6) ; the energy of the particle E > U . U(x) Fig. 6 Solution. Throughout the region x > 0, the wave function has the form (25.1), while in the region x < its form is (25.2). The constants A and B are determined from the condi- tion that i/i and dift/dx are continuous at * = : 1+B=A, k x {\-B) =M, whence A = 2k x i{k x +kt\ B = {k x -k 2 )l{k x +k 2 ). The reflection coefficient f is (25.4) R = / v-^y = / p ± -p 2 \* \k x +kj Kpx+pi)' For E = U (k 3 = 0), R becomes unity, while for E -> oo it tends to zero as (C7 /4J?) 2 . f In the limiting case of classical mechanics, the reflection coefficient must become zero. The expression obtained here, however, does not contain the quantum constant at all. This apparent contradiction is explained as follows. The classical limiting case is that in which the de Broglie wavelength of the particle A ~ hip is small in comparison with the characteristic dimensions of the problem, i.e. the distances over which the field U(x) changes noticeably. In the schematic example considered, however, this distance is zero (at the point x = 0), so that the passage to the limit cannot be effected. 78 Schrddinger' s Equation §25 Problem 2. Determine the transmission coefficient for a rectangular potential barrier (Fig. 7). U„ y(x) Fig. 7 Solution. Let E be greater than U , and suppose that the incident particle is moving from left to right. Then we have for the wave function in the different regions expressions of the form for * < 0, iff = e ik i x +Ae- ik i x , for0<x<a,ip= Be ik * x +B'e- ik i x , for * > a, ifi = Ce ik i x (on the side x > a there can be only the transmitted wave, propagated in the positive direc- tion of x). The constants A, B, B' and C are determined from the conditions of continuity of tfi and dtfi/dx at the points x = and a. The transmission coefficient is determined as D = ^|C| 2 /^i = |C| 2 . On calculating this, we obtain D = lf\i f\a For E < U , &a is a purely imaginary quantity; the corresponding expression for D is obtained by replacing k% by ik 2 , where Hk z = \/[2m(U — E)]: D = lantity ; = V[2m( t/Jj /c 2 (&! 2 +/c 2 2 ) 2 sinh«fl#c 8 +4ife 1 a f2 a ' Problem 3. Determine the reflection coefficient for a potential wall defined by the formula U(x) = J7 /(l +e~ ax ) (Fig. 5); the energy of the particle is E > U . Solution. Schrodinger's equation is d 2 if/ 2m/ d 2 «/r 2m/ U \ —-+—[E °—U = d* 2 h 2 \ l+e-^J We have to find a solution which, as x -> +oo, has the form iff = constant x e ik * x . We introduce a new variable £ = — e~ aX (which takes values from -co to 0), and seek a solution of the form = £-**,/««;(£), where w(£) tends to a constant as £ -> (i.e., as x ->■ co). For w(£) we find an equation of hypergeometric type: £(l-£) w "+(l-2ikjK)(l-0w'^k 2 *-k 1 *)wl«? = 0, §25 The transmission coefficient 79 which has as its solution the hypergeometric function w = F(i[^ 1 -^ 2 ]/a,-^ 1 +A 2 ]/a,-2^ 2 /a+l, $) (we omit a constant factor). As £ ->■ 0, this function tends to 1, i.e. it satisfies the condition imposed. The asymptotic form of the function ^ as £ -*■ — oo (i.e. x -» — oo) isf if, « |-^,/«[C 1 (-|) i (V* ; i)/»+C 2 (-^ fc i+ A: i >/ a ] = (-l)-**./°[C 1 e < *i*+C a e-«*i a! ], where r(-2* 1 /a)r(-2tfe 2 /a+ 1) C 1 = c, = r(_z-(^ 1 +^)/a)r(-^ 1 +A 2 )/a+l) T(2ife 1 /a)r(-2«fe 2 /a+l) r(^ 1 -^ 2 )/a)r(t(^-^ 2 )/a+ 1)* The required reflection coefficient is R = IC2/CJ 2 ; on calculating it by means of the well- known formulae we have T(x+1) =xT(x), r(*)r(l-*) =7r/sin7r^, r = / sinhfrrfo-fej/tt] \ 2 Vsinh[7r(^ 1 +^ 2 )/a] / For E = U (& 2 = 0), i? becomes unity, while for E -> co it tends to zero as 7rU n \ 2 2m /7rU \ z 2m e -i*\/(2mE)/ah, In the limiting case of classical mechanics, R becomes zero, as it should. Problem 4. Determine the transmission coefficient for a potential barrier denned by the formula U(x) = U Q lcosh 2 ctx (Fig. 8) ; the energy of the particle is E < U Q . U(x) Fig. 8 Solution. The Schrodinger's equation is the same as that obtained in the solution of Problem 4, §23 ; it is necessary merely to alter the sign of Uo and to regard the energy E now as positive. A similar calculation gives the solution 4> = (l~P)- ik ' 2 "F[-iklx-s, -ik/x+s+1, -ik/a+ !,&!-£)], (1) f See formula (e.6), in each of whose two terms we must take only the first term of the expansion i.e. replace the hypergeometric functions of 1/a by unity. 80 Schrddinger' s Equation §25 where $ = tanh cox, k = y/{lmE)\%, SmU - -i-»A<-m This solution satisfies the condition that, as x -*■ oo (i.e. as $ -> 1, (1 — f) *» 2e~ x ), the wave function should include only the transmitted wave (~e* fcz ). The asymptotic form of the wave function as x -> — oo (£ -*■ ~ 1) is found by transforming the hypergeometric function with the aid of formula (e.7) : rmm - ik) rr - ik)va - ik) j, „ e -ik x v ' v L +e ikx 1 L± i . (2) T(-s)T(l+s) T(-ik-s)r(-ik+s+l) K J Talcing the squared modulus of the ratio of coefficients in this function, we obtain the follow- ing expression for the transmission coefficient D = 1 — R : sinh 2 (7r£/a) (if 8mt/o/£ 2 <x 2 <l), or D = sinh 2 (7rfc/a) + cos 2 [£tt V( 1 - 8m U /h 2 oi. 2 )] sinh 2 (7r£/a) sinh 2 (7r£/a)+cosh 2 |>- V(8m£y# 2 a 2 - 1)] (if 8tnUo/h 2 ofi > 1). The first of these formulae holds also for the case Uo < 0, i.e. when the particle is passing over a potential well instead of a potential barrier. It is interesting to note that in that case D = 1 if 1 +8m\Uo\lfr 2 a? = (2« + l) 2 ; thus, for certain values of the depth | Uq\ of the well, particles passing over it are not reflected. This is evident from equation (2), where the term in e~ ikx vanishes for positive integral s. CHAPTER IV ANGULAR MOMENTUM §26. Angular momentum In §15, to derive the law of conservation of momentum, we have made use of the homogeneity of space relative to a closed system of particles. Besides its homogeneity, space has also the property of isotropy: all directions in it are equivalent. Hence the Hamiltonian of a closed system cannot change when the system rotates as a whole through an arbitrary angle about an arbitrary axis. It is sufficient, as in §15, to require the fulfilment of this con- dition for an infinitely small rotation. Let 8<p be the vector of an infinitely small rotation, equal in magnitude to the angle 8<f> of the rotation and directed along the axis about which the rotation takes place. The changes Sr a (in the radius vectors r a of the par- ticles) in such a rotation are well known to be 8r = 8<p x r a . An arbitrary function j/rfo, r 2 , ... ) is thereby transformed into the function ^1+8^,^+8^,...) = «A(r 1 ,r 2 ,...)+S8r . V a ^ = #*i» r 2 , ... )+ S 8<p x r a . V a ^ = (l+8<p.Sr a xV a )0(r 1 ,r 2 ,...). The expression l+8<p.Sr a xV« can be regarded as the "operator of an infinitely small rotation". The fact that an infinitely small rotation does not alter the Hamiltonian of the system is expressed (see §15) by the commutability of the "rotation operator" with the operator i?. Since the operator of multiplication by unity commutes with any operator, while 8<p is a constant vector, this condition reduces to the relation (2 r a x V«)#-#(2 r a x V a ) = 0, (26.1) which expresses a certain law of conservation. The quantity whose conservation for a closed system follows from the property of isotropy of space is the angular momentum of the system. Thus the operator Sr a x V a must correspond exactly, apart from a constant factor, to the total angular momentum of the system, and each of the terms r a x V a of this sum corresponds to the angular momentum of an individual particle. 81 82 Angular Momentum §26 The coefficient of proportionality must be put equal to — ih\ this follows immediately, because then the expression for the angular momentum operator of a particle is —ihrx V = rxp and corresponds exactly to the familiar classical expression rxp. Henceforward we shall always use the angular momentum measured in units of h. The angular momentum operator of a particle, so defined, will be denoted by 1, and that of the whole system by L. Thus we have for the angular momentum component operators of a particle the expressions ht x = yp z —zpy, ht y = zp x —xp z , hl z = xp y —yp x . (26.2) For a system which is in an external field, the angular momentum is in general not conserved. However, it may still be conserved if the field has a certain symmetry. Thus, if the system is in a centrally symmetric field, all directions in space at the centre are equivalent, and hence the angular momen- tum about this centre will be conserved. Similarly, in an axially symmetric field, the component of angular momentum along the axis of symmetry is conserved. All these conservation laws holding in classical mechanics are valid in quantum mechanics also. In a system where angular momentum is not conserved, it does not have definite values in the stationary states. In such cases the mean value of the angular momentum in a given stationary state is sometimes of interest. It is easily seen that, in any non-degenerate stationary state, the mean value of the angular momentum is zero. For, when the sign of the time is changed, the energy does not alter, and, since only one stationary state corresponds to a given energy level, it follows that when t is changed into — t the state of the system must remain the same. This means that the mean values of all quantities, and in particular that of the angular momentum, must remain unchanged. But when the sign of the time is changed, so is that of the angular momentum, and we have L = — L, whence it follows that L = 0. The same result can be obtained by starting from the mathematical definition of the mean value L as being the integral of i/i*Lift. The wave functions of non- degenerate states are real (see the end of §18). Hence the expression L = — ih J 0(Sr o x V o )0d# is purely imaginary, and since L must, of course, be real, it is evident that L=0. Let us derive the rules for commutation of the angular momentum operators with those of co-ordinates and linear momenta. By means of the relations (16.2) we easily find {! x ,x} = 0, {l x ,y} = tar, {t x ,z} = —iy, tfv»3'}= » &»*}=**» {/„*}= —tar, \ (26.3) {/„*} = 0, {t z ,x} = iy, {t z ,y} = -ix. §26 Angular momentum 83 For instance, Ly-yL = (m)(yh-4v)y-y(yfa-4v)(W) = -(zlh){$v,y} = iz- All the relations (26.3) can be written in tensor form as follows: {h,x k } =te m x u (26.4) where e iU is the antisymmetric unit tensor of rank three,f and summation is implied over those suffixes which appear twice (called dummy suffixes). It is easily seen that a similar commutation rule holds for the angular momentum and linear momentum operators : {4 h) = i*ikipi- (26.5) By means of these formulae, it is easy to find the rules for commutation of the operators ! x , t y , l z with one another. We have HUv-tJx) = L(zfix- x Pz)-(zpx-x$ g )L = (lxZ-zL)Px—x(t x } z —$J x ) = —iyfrx+ixpy = iht z . {l v Jz}=iL {tzJx}=il v , {/«,/,}=«/„ (26.6) Thus or {4,4} =**«,/,. (26.7) Exactly the same relations hold for the operators L x ,L y ,L z of the total angular momentum of the system. For, since the angular momentum oper- ators of different individual particles commute, we have, for instance, 5'«»|''«— S/asS/ay = 2>{hvhz—hzhv) = »S t ax . Thus {t vt l z }=iL x , {L z ,l x } =it y , {L x ,L y }=iL z . (26.8) The relations (26.8) show that the three components of the angular momen- tum cannot simultaneously have definite values (except in the case where all three components simultaneously vanish: see below). In this respect the angular momentum is fundamentally different from the linear momentum, whose three components can simultaneously have definite values. f The antisymmetric unit tensor of rank three, e m (also called the unit axial tensor), is denned as a tensor antisymmetric in all three suffixes, with e nz = 1 . It is evident that, of its 27 components, only 6 are not zero, namely those in which the suffixes i, k, I form some permutation of 1, 2, 3. Such a com- ponent is +1 if the permutation i, k, I is obtained from 1, 2, 3 by an even number of transpositions ot pairs of figures, and is - 1 if the number of transpositions is odd. Clearly e m e ikm = 2Sim emenci = 6. The components of the vector C = AxB which is the vector product of the two vectors A and B can be written by means of the tensor em in the form c, = 84 Angular Momentum §26 From the operators L x , L yy L z we can form the operator L x 2 +L y 2 +L*, which can be regarded as the operator of the square of the modulus of the angular momentum vector, and which we denote by L 2 : L* = Lf+Lf+L*. (26.9) This operator commutes with each of the operators L x ,L yi L e : {L\t x }=0, {L*,Z„}=0, {L*,4}=0. (26.10) Using (26.8), we have {l x \l z } =l x {l x ,l z }+{L x ,t z }l x = l^Lgljy + LyLgc), {Ly t Lg} = l^Lgljy + LyL;,)), {L,\L.} =o. Adding these equations, we have {L 2 , L z ) = 0. Physically, the relations (26.10) mean that the square of the angular momentum, i.e. its modulus, can have a definite value at the same time as one of its components. Instead of the operators L x , L y it is often more convenient to use the complex combinations L+ = Lx+iLy, ■£- = L x —iLy. (26.11) It is easily verified by direct calculation using (26.8) that the following commutation rules hold : {L + , I-} = 2t z , {L z , L + } =L + , {l z ,l-}=-l- t and it is also not difficult to see that L2 = t + t-+t z *-l z = L-L + +L z *+l*. (26.13) Finally, we shall give some frequently used expressions for the angular momentum operator of a single particle in spherical polar co-ordinates. Defining the latter by means of the usual relations * = r sin 8 cos <f>, y = r sin 6 sin <f>, z = r cos 0, we have after a simple calculation U = -A (26.14) /. = e±**[ + — +* cot 9— ). (26.15) §27 Eigenvalues of the angular momentum 85 Substitution in (26.13) gives the squared angular momentum operator of the particle : 12 r 1 ^ 1 d f d \i It should be noticed that this is, apart from a factor, the angular part of the Laplacian operator. §27. Eigenvalues of the angular momentum In order to determine the eigenvalues of the component, in some direction, of the angular momentum of a particle, it is convenient to use the expression for its operator in spherical polar co-ordinates, taking the direction in question as the polar axis. According to formula (26.14), the equation ty = 1$ can be written in the form -i ty/ty = Irf. (27.1) Its solution is *!> = /(r,0)A* where/(r, 0) is an arbitrary function of r and 0. If the function $ is to be single- valued, it must be periodic in <f>, with period 2tt. Hence we findf h = m, where m = 0, ±1,± 2, ... . (27.2) Thus the eigenvalues l z are the positive and negative integers, including zero. The factor depending on <j>, which characterises the eigenfunctions of the operator t„ is denoted by Q n (t) = (27r)-W»* (27.3) These functions are normalised so that j ® m *(<t>)® m <<!>) d0 = h mm '. (27.4) The eigenvalues of the ^-component of the total angular momentum of the system are evidently also equal to the positive and negative integers: L s = M, where M = 0,±1, ±2, ... (27.5) (this follows at once from the fact that the operator L z is equal to the sum of the commuting operators 4 for the individual particles). Since the direction of the ^-axis is in no way distinctive, it is clear that the same result is obtained for L x , i y and in general for the component of the angular momentum in any direction: they can all take integral values only. At first sight this result may appear paradoxical, particularly if we apply it to two directi ons infinitely close to each other. In fact, however, it must i f J h V US l° mary no r tation . fo , r the eigenvalues of the angular momentum component is m, which also denotes the mass of a particle, but this should not lead to any confusion. 86 Angular Momentum §27 be remembered that the only common eigenfunction of the operators L x , L y , L z corresponds to the simultaneous values L x = L y = L z = u; in this case the angular momentum vector is zero, and consequently so is its projection upon any direction. If even one of the eigenvalues L x ,L y ,L s is not zero, the operators L x ,L y ,L e have no common eigenfunctions. In other words, there is no state in which two or three of the angular momentum components in different directions simultaneously have definite values differ- ent from zero, so that we can say only that one of them is integral. The stationary states of a system which differ only in the value of L z have the same energy; this follows from general considerations, based on the fact that the direction of the sr-axis is in no way distinctive. Thus the energy levels of a system whose angular momentum is conserved (and is not zero) are always degenerate, f Let us now look for the eigenvalues of the square L 2 of the angular momen- tum. We shall show how these values may be found, starting from the commutation conditions (26.8) only. We denote by \Jj m the wave functions of the stationary states belonging to one degenerate energy level and distin- guished by the value of L z = M. Besides the energy, the square L 2 of the angular momentum also has a (single) definite value in these states. J First of all, we note that since the difference L 2 -4 2 = £* 2 + V is equal to the operator of the essentially positive physical quantity L x 2 +L y 2 , it follows that, for a given value of the squared angular momentum L 2 and any possible eigenvalue of the quantity L s , the inequality L 2 ^ L s 2 , or - VL 2 < L z < + VL 2 , (27.6) must hold. Thus the possible values of L z (for a given L 2 ) are bounded by certain upper and lower limits; we denote by L the integer corresponding to the greatest value of \L Z \. Next, by applying the operator L Z L ± to the eigenfunction j/tm of the operator L z and using the commutation rule {L z , £ ± } = ±L ± (26.12), we obtain LzLJjM = {M± l)£ ± 0a*. t This is a particular case of the general theorem, mentioned in §10, which states that the levels are degenerate when two or more conserved quantities exist whose operators do not commute. Here the components of the angular momentum are such quantities. t Here it is supposed that there is no additional degeneracy leading to the same value of the energy for different values of the squared angular momentum. This is true for a discrete spectrum (except for the case of what is called accidental degeneracy in a Coulomb field; see §36) and in general untrue for the energy levels of a continuous spectrum. However, even when such additional degeneracy is present, we can always choose the eigenfunctions so that they correspond to states with definite values of L 2 , and then we can choose from these the states with the same values of E and L 2 . This is mathe- matically expressed by the fact that the matrices of commuting operators can always be simultaneously brought into diagonal form. In what follows we shall, in such cases, speak, for the sake of brevity, as if there were no additional degeneracy, bearing in mind that the results obtained do not in fact depend on this assumption, by what we have just said. §27 Eigenvalues of the angular momentum 87 Hence we see that the function Lj, M is (apart from a normalisation constant) the eigenfunction corresponding to the value M± 1 of the quantity L z - we can write ' fa+i = constant xL + fa, ^m-i = constant x Z_0 M . (27.7) If we put M = L in the first of these equations, we must have identically £-4l = 0, (27.8) since there is by definition no state M > L. Applying the operator L- to this equation and using the relation (26.13), we obtain LXrf L = (Ifi-lf-L^ = o. Since, however, the fa are common eigenfunctions of the operators L 2 and L z , we have so that the equation found above gives L2=Z(£+1). (279) Formula (27.9) determines the required eigenvalues of the square of the angular momentum; the number L takes all positive integral values, including zero. For a given value of Z, the component L z = M of the angular momen- tum can take the values M = L,L-1,...,-L, (27.10) i.e. 2L + 1 different values in all. The energy level corresponding to the angular momentumf L thus has (2L + l)-fold degeneracy. A state with angular momentum L = (when all three components are also zero) is not degenerate; we notice that the wave function of such a state is spherically symmetric. This follows from the mere fact that, when acted on by the angular momentum operator, it becomes zero, i.e. it is unchanged as a result of any infinitely small rotation. For the angular momentum of a single particle we write formula (27.9) in the form I 2 ='('+!)> (27.11) i.e. we denote the angular momentum of an individual particle by the small letter /. J Let us calculate the matrix elements of the quantities L x and L y in a representation in which L z and V, as well as the energy, are diagonal (M. Born, W. Heisenberg and P. Jordan 1926). First of all, we note that, since the operators L x and L y commute with the operator 8, their matrices are diagonal with respect to the energy, i.e. all matrix elements for transitions t We shall often, for the sake of brevity, and in accordance with custom, speak of the "angular momentum L of a system, understanding by this a momentum whose greatest possible component 88 Angular Momentum §28 between states of different energy (and different angular momentum L) are zero. Thus it is sufficient to consider the matrix elements for transitions within a group of states with different values of M, corresponding to a single degenerate energy level. It is seen from formulae (27.7) that, in the matrices of the operators L+ and L-, only those elements are different from zero which correspond to transitions M+ 1 -> M and M— 1 -> M respectively. Taking this into account, we find the diagonal matrix elements on both sides of the equation (26.13), obtainingf L(L+1) = (L+) m , M -i(L-)m-i,m+M*-M. Noticing that, since the operators L x and L y are Hermitian, (L-)m-1,M = (L+)*m,m-i> we can rewrite this equation in the form \(L + )m,m-i\ 2 = L(L+1)-M(M-1) = (L-M+1)(L+M), whencej (L+)m,m-i = (L-)m-i,m = V[(L+M)(L-M+1)]. (27.12) Hence we have for the non-zero matrix elements of the quantities L x and L y themselves (L x )m,m-i = (L x ) M -i,m = * V[(L+M)(L-M+ 1)], {L v )m,m-x = -(L y ) M -i,M = -VV[(L+M)(L-M+1)]. In the corresponding formulae for the angular momentum of a particle, we must write /, m instead of L, M. §28. Eigenfunctions of the angular momentum The wave function of a particle is not completely determined when the values of l 2 and l z are prescribed. This is seen from the fact that the expres- sions for the operators of these quantities in spherical polar co-ordinates contain only the angles 6 and <f>, so that their eigenfunctions can contain an arbitrary factor depending on r. We shall here consider only the angular part of the wave function which characterises the eigenfunctions of the f In the symbols for the matrix elements, we omit for brevity all suffixes with respect to which they are diagonal (including L). % The choice of sign in this formula corresponds to the choice of the phase factors in the eigenfunc- tions of the angular momentum. § 28 Eigenfunctions of the angular momentum 89 angular momentum, and denote this by Y^fi, with the normalisation condition J|rj»do«i, where do = sin 6 ddd<f> is an element of solid angle. We shall see that the problem of determining the common eigenfunctions of the operators i 2 and l s admits of separation of the variables B and <f> and these functions can be sought in the form Ylm = «Ufl©,„(0), (28.1) where O m (<f>) are the eigenfunctions of the operator / 3 , which are given by formula (27.3). Since the functions O m are already normalised by the condi- tion (27.4), the 0, m must be normalised by the condition V J|0, w | 2 sin0d0 = l. ( 28.2) o The functions Y lm with different / or m are automatically orthogonal: / / Y Vm * Yim sin ddd<f> = M..', (28.3) o o as being the eigenfunctions of angular momentum operators corresponding to different eigenvalues. The functions OJfl separately are themselves orthogonal (see (27.4)), as being the eigenfunctions of the operator l e cor- responding to different eigenvalues m of this operator. The functions (0) are not themselves eigenfunctions of any of the angular momentum operators • they are mutually orthogonal for different /, but not for different m. The most direct method of calculating the required functions is by directly solving the problem of finding the eigenfunctions of the operator I 2 written in spherical polar co-ordinates (formula (26.16)). The equation f 2 <£ = IV is Substituting in this equation the form (28.1) for 0, we obtain for the function &i m the equation 1 d /. d0 Zw \ m 2 5n7SV Sm '^rJ-^ - +/ ( /+1 ) - - °- (28.4) This equation is well known in the theory of spherical harmonics. It has solutions satisfying the conditions of finiteness and single-valuedness for positive mtegral values of / > | m |, in agreement with the eigenvalues of the angular momentum obtained above by the matrix method. The correspond- ing solutions are what are called associated Legendre polynomials P?(cos 6) 90 Angular Momentum §28 (see §c of the Mathematical Appendices). Using the normalisation condition (28.2), we findf €U0) = {-l) m iWmi+W-^)W+m)\]Pi m (cos9). (28.5) Here it is supposed that m ^ 0. For negative m, we use the definition ©i.-i»i=(-1)"®«i«i- ( 28 - 6 ) In other words, @i m for m < is given by (28.5) with \m\ instead of m and the factor (— l) m omitted. For m = 0, the associated Legendre polynomials are called simply Legendre polynomials Pi(cos 6) ; we have 0*o = i l V[«2/+ l)]P,(oo8 0). (28.7) Thus the eigenfunctions of the angular momentum are mathematically just spherical harmonic functions normalised in a particular way. From (27.3) and (28.6) it is seen that functions differing in the sign of m are related by the equation (-l) l - m Y l ,- m = Y lm *. (28.8) We shall make some remarks concerning the eigenfunctions of the angular momentum. For / = (so that m = also) this function reduces to a con- stant. In other words, the wave functions of the states of a particle with zero angular momentum depend only on r, i.e. they have complete spherical symmetry. For a given m, the values of / starting from \m\ denumerate the successive eigenvalues of the quantity / in order of increasing magnitude. Hence, from the general theory of the zeros of eigenfunctions ( §21), we can deduce that the function @ lm becomes zero for l—\m\ different values of the angle 6; in other words, it has as nodal lines /— \m\ "lines of latitude" on the sphere. If the complete angular functions are taken with the real factors cos m<j> or sin m<f> instead of J «±*»#, they have as further nodal lines \m\ "lines of longi- tude"; the total number of nodal lines is thus /. Finally, we shall show how the functions S lm may be calculated by the matrix method. This is done similarly to the calculation of the wave func- tions of an oscillator in §23. We start from the equation (27.8) : t+Y a = 0. Using the expression (26.15) for the operator /+ and substituting Yu = (27r)-V 7 *0 w (0), we obtain for ®u the equation d0„/d0-Zcot0 0„ =0, t The choice of the phase factor is not, of course, determined by the normalisation condition. The definition (28.5) used in this book is the most natural from the viewpoint of the theory of addition of angular momenta (see Chapter XIV). It differs by a factor i l from the one usually adopted. X Each such function corresponds to a state in which l z does not have a definite value, but can have the values ±m with equal probability. §29 Matrix elements of vectors 91 whence ®u = constant X sin'0. Determining the constant from the normali- sation condition, we find % = (-*M£(2/+1)!]2-'(1/Z!) sin'0. (28.9) Next, using (27.12), we write *-^,w+l = ('-)m,m+l^im = V[(l-m)(l+m+l)]Y ln . A repeated application of this formula gives V[(l-m)\l(l+m)\]Y lm = [(20!]- 1 /2(/_)^F„. The right-hand side of this equation is easily calculated by means of the expression (26.15) for the operator L. We have L[f(ey™<t>] = e^m-m & { n i-me d(/sin^)/d(cos d). A repeated application of this formula gives {!_)l-m e -04>® n = fdrrul, ^-mg d *-m( sin * q 0^)/ d ( cos Qf-m^ Finally, using these relations and the expression (28.9) for ZZ , we obtain the formula which is the same as (28.5). §29. Matrix elements of vectors Let us again consider a closed system of particles ;f let/ be any scalar physical quantity characterising the system, and /the operator corresponding to this quantity. Every scalar is invariant with respect to rotation of the co-ordinate system. Hence the scalar operator / does not vary when acted on by a rotation operator, i.e. it commutes with a rotation operator. We know, however, that the operator of an infinitely small rotation is the same, apart from a constant factor, as the angular momentum operator, so that {/, L.) = {/, Lj = {/ 4} = (29.1) (and also {/, L 2 } = 0). From the commutability of / with the angular momentum operator it follows that, in a representation where L 2 and L z are diagonal, the matrix of the quantity / will also be diagonal with respect to the suffixes LM. We shall conventionally denote by n all the remaining suffixes which define the t AH the results in this section are valid also for a particle in a centrally symmetric field (and in general whenever the total angular momentum of the system is conserved). 92 Angular Momentum §29 state of the system, and we shall show that the matrix elements f^L are independent of the suffix M. To do this, we use the commutability of /with /£+-£+/= 0. Let us write down the matrix element of this equation corresponding to the transition «, L, M -> «', L, M— 1. Taking into account the fact that the matrix of the quantity L + has only elements with n,L,M-+ n, L, M—l, we obtain nLM n'LM nLM n L,M-l J n'LM{^+)n'L,M-l— \L>+)nL,M-lJ n'L,M-l = «, and since the matrix elements of the quantity L+ are independent of the suffix n, we find nLM nL,M-l J n'LM =Jn'L,M-l, whence it follows that all the quantities ffilh for different M (the other suffixes being the same) are equal. Thus the matrix elements of the quantity / that are different from zero will be nLM nL • J n'LM =Jn'L, (29.2) where ffi L denotes quantities depending on the values of the suffixes «, «', L. If we apply this result to the Hamiltonian itself, we obtain our previous result that the energy of the stationary states is independent of M, i.e. that the energy levels have (2L+l)-fold degeneracy. Next, let A be some vector physical quantity characterising a closed system. When the system of co-ordinates is rotated (and, in particular, when the operator of an infinitely small rotation, i.e. the angular momentum opera- tor, is applied), the components of a vector are transformed into linear functions of one another. Hence, as a result of the commutation of the operators L t with the operators A t> we must again obtain components of the same vector, A%. The exact form can be found by noticing that, in the particular case where A is the radius vector of the particle, the formulae (26.3) must be obtained. Thus we find the commutation rules {l x ,A x } = 0, {l x ,A y } = iA z , {L X ,A Z }= -iA y , ... (29.3) (the remaining rules are obtained by cyclic permutation of the suffixes x, y, z), or {l u A k } = ie m A x . (29.4) These relations enable us to obtain several results concerning the form of the matrices of the components of the vector A (M. Born, W. Heisenberg and P. Jordan 1926). First of all, it is possible to derive selection rules which determine the transitions for which the matrix elements can be different from zero. We shall not go through the fairly lengthy calculations here, however, since it will appear later (§107) that these rules are actually a direct §29 Matrix elements of vectors 93 consequence of the general transformation properties of vector quantities and can be derived from the latter with hardly any calculation at all. Here we shall merely give the rules, without proof. The matrix elements of all the components of a vector can be different from zero only for transitions in which the angular momentum L changes by not more than one unit : L-+L or L±l. (29.5) There is a further selection rule which forbids transitions between any two states with L = 0. This rule is an obvious consequence of the complete spherical symmetry of states with angular momentum zero. The selection rules for the angular momentum component M are different for the different components of a vector: the matrix elements can be different from zero for transitions where M changes as follows: for A+ = A x +iA y , M-+M-1, n for A- = A x -iA y , M-+M+1, (29.6) for A z , M^M. J Moreover, it is possible to determine a general form for the matrix elements of a vector as functions of the number M. These important and frequently used formulae are given here, also without proof, since they are actually a particular case of more general relations derived in §107 for any tensor quantities. The non-zero matrix elements of the quantity A z are given by the formulae M (A^ n ^ M = A^ L Wn'LM V[L(L+1)(2L+1) fn'L> W^ 1 . M = y £(2£ _ 1)(2£+1) ^ .,- 1 . | (29.7) yL 2 —M 2 L(2L-1)(2L+1) nL Here the A^ L . are quantities! independent of M and related by the "Her- mitian" condition a I>l> = (^£?>. (29.8) which follows directly from the fact that the operator A z is Hermitian. t These quantities are sometimes called reduced matrix elements. The appearance in formulae (29.7) and (29.9) of denominators which depend on L is in accordance with the general notation used in §107. The convenience of these denominators is shown, in particular, by the simple form of equation (29.12) for the matrix elements of the scalar product of two vectors. It may be noted for reference that for the vector L itself the reduced matrix elements are L\ = VW+l)(2L+l)l L^ 1 = Ll_ x = 0, since the matrix of L z is diagonal with respect to L and the diagonal elements are equal to M. The same values are, of course, obtained from a comparison of formulae (29.9) and (27.12). 94 Angular Momentum §29 The matrix elements of the quantities A- and A + are also determined by the A^\j. The non-zero matrix elements of A- are (A \nL ,M-1 _ 12 /v ' An L Wwlm -J L(L+l)(2L+ir n ' L ' (A \nL,M-l _ ± '2 L/jnL 1 -V,l-i,m ./ L{2 L-1)(2L+1) n '' L ~ v (A )n ,L-l,M-l - _ ^l^Zl^l^l A n,L-l (29.9) (L + M-1)(L + M) L(2L-1)(2L+1) The matrix elements of A+ need not be written out separately: since A x and A y are real we have, using (11.9), (4)ZI&M> = U-)nLM M ']*- (29.10) It is useful to notice a formula which expresses the matrix elements of a scalar A . B (where A, B are two vector physical quantities) in terms of the coefficients A^, B n J- L , in formulae (29.7)-(29.9); in applications such products are usually involved. The calculation is conveniently carried out by writing the operator A . B in the form A . B = i(A + B-+A~B+)+A z B z . (29.11) It is evident a priori that the matrix of A . B (like that of any scalar) is diagonal with respect to L and M. A calculation by means of formulae (29.7)-(29.9) gives the result (A.B)»f* f = i ^- r 2^-i" B »"'f. (29.12) n",L" where L" takes the values L, L ± 1. PROBLEMS Problem 1. Determine the matrix elements (with respect to the eigenfunctions of the angular momentum) of a unit vector n along the radius vector of the particle. Solution. The matrix elements of a polar vector for an individual particle are non-zero only for transitions / -*■ l± 1 (see §30). Their dependence on the quantum number m is given by the general formulae (29.7)-(29.9), so that it is sufficient to calculate the coefficient n \-i = ( n l X )* (corresponding to A % _ t in those formulae). To do this, we can find, for example, the matrix element of n z = cos 6 as follows: 7 1 n (^I 1)(2 / + i) "J" 1 = (C ° S d) "~'° = / 0; - 1 ' * cos 6 • ® 10 sin 6 dd > with the functions ©jo given by (28.7); the integral is calculated from formula (107.15) with h = \,h = 1—1. The result is Problem 2. Average the tensor nitik—^ik (where n is a unit vector along the radius vector of a particle) over a state where the magnitude but not the direction of the vector 1 is given (i.e. l z is indeterminate). §30 Parity of a state 95 Solution.^ The required mean value is an operator which can be expressed in terms of the operator 1 alone. We seek it in the form nink ~ %$ik — a[hh + tJi — §Sifc/(/+ 1)] ; this is the most general symmetrical tensor of rank two with zero trace that can be formed from the components of I. ^ To determine the constant a we multiply this equation on the left by U and on the right by h (summing over i and k). Since ^the vector n is perpendicular to the vector M = rxp, we have mU = 0. The product UlihU = (l 2 ) 2 is replaced by its eigenvalue l\l+\) z , and the product UUUh is transformed by means of the commutation relations (26.7) as follows: hhhh = hhhh—iemhhh = fi 2 ) 2 -¥ e mk(tih-hli) = (l 2 ) 2 +ieikieikmUm = (12)2 _ J2 = / 2 (/+l) 2 -/(/+l) (using the fact that eme m ki = 2Sj m ). After a simple reduction we obtain the result a= -l/(2/-l)(2/+3). §30. Parity of a state Besides the parallel displacement of the co-ordinate system (used in §15) and the rotation of it (used in §26), there is another transformation which leaves unaltered the Hamiltonian of a closed system, f This is what is called the inversion transformation, which consists in simultaneously changing the sign of all the co-ordinates. In classical mechanics, the invariance of Hamilton's function with respect to inversion does not lead to a conser- vation law, but the situation is different in quantum mechanics. Let us denote by / the inversion operator; its effect on a function is to change the sign of all the co-ordinates. The invariance of ff with respect to inversion means that fil-lfi = 0. (30.1) The operator / also commutes with the angular momentum operators : {/, L x } = {/, L y ) = {/, L z } = o, {/, L*} = o (30.2) (on inversion, both the co-ordinates themselves and the operators of differ- entiation with respect to them change sign, so that the angular momentum operators remain unchanged). It is easy to find the eigenvalues J of the inversion operator, which are deter- mined by the equation U = fy. f The same is true of a system in a centrally symmetric field. 96 Angular Momentum §30 To do this, we notice that a double application of the operator / amounts to identity : no co-ordinate is altered. In other words, we have Piff = Pift — fa i.e. P = 1, whence / = ±1. (30.3) Thus the eigenfunctions of the inversion operator are either unchanged or change in sign when acted upon by this operator. In the first case, the wave function (and the corresponding state) is said to be even, and in the second it is said to be odd. The equation (30.1) thus expresses the "law of conservation of parity"; if the state of a closed system has a given parity (i.e. if it is even, or odd), then this parity is conserved. The physical meaning of equations (30.2) is that the system can have defin- ite values of L and M and, at the same time, a definite parity of its state. We can also say that all states differing only in the value of M have the same parity. This can be shown by starting from the relation L4-1U = o and proceeding in exactly the same way as in obtaining the result (29.2). When the inversion transformation is applied to scalar quantities, either they do not change at all (true scalars) or they change sign (what are called pseudoscalars).-\ If a physical quantity / is a true scalar, its operator com- mutes with /: 7/-// = 0. (30.4) It follows from this that, if the matrix of I is diagonal, then the matrix of / is diagonal with respect to the suffix which shows the parity of the state, i.e. only the matrix elements for transitions u -> u and g ->• g are not zero (the suffixes g and u denote even and odd states respectively). For the operator of a pseudoscalar quantity we have // = — //, or //+// = 0; (30.5) / anticommutes with /. The matrix element of this equation for a transition and since I gg =1 we have f gg = (we omit all suffixes apart from that showing the parity). Similarly we find that f uu — 0. Thus, in the matrix of a pseudoscalar quantity, only those elements can be different from zero which are non-diagonal with respect to the parity suffix (transitions with change of parity). \ An example of a pseudoscalar is the scalar product of an axial and a polar vector. §31 Addition of angular momenta 97 Similar results are obtained for vector quantities. The operators of a polar vectorf anticommute with /, and in their matrices (in a representation where 1 is diagonal) only the elements for transitions with change of parity are not zero. The operators of an axial vector, however, commute with / , and their matrices have non-zero elements only for transitions without change of parity. It is useful to point out another method of obtaining these results. For example, the matrix element of a scalar / for a transition between states of opposite parity is the integral f ug = $ tft u *ftfi g dq, where the function \\t g is even and tj/ u is odd. When all the co-ordinates change sign, the integrand does so if / is a true scalar; on the other hand, the integral taken over all space cannot change when the variables of integration are re-named. Hence it follows that/ u? = -/ ua , i.e. / w s 0. Let us determine the parity of the state of a single particle with angular momentum /. The inversion transformation (x -> — x, y -> —y, z -> —z) is, in spherical polar co-ordinates, the transformation r^r, 6-+TT-9, ^->tt+0. (30.6) The dependence of the wave function of the particle on the angle is given by the eigenfunction Y Jm of the angular momentum, which, apart from a constant which is here unimportant, has the form P, m (cos &)e im *. When <j> is replaced by ir+<f>, the factor e im * is multiplied by ( — l) m , and when 6 is replaced by 77—0, P, m (cos0) becomes P^-cos 0) = (-l) l - m P J wl (cos 0). Thus the whole function is multiplied by ( — l) 1 (independent of m, in agreement with what was said above), i.e. the parity of a state with a given value of / is / = (-1) 1 . (30.7) We see that all states with even I are even, and all those with odd / are odd. A vector physical quantity relating to an individual particle can have non- zero matrix elements only for transitions with / -> / or /±1 (§29). Remem- bering this, and comparing formula (30.7) with what was said above regarding the change of parity in the matrix elements of vectors, we reach the result that the matrix elements of a polar vector are non-zero only for transitions with / -> 7±1, and those of an axial vector for transitions with / -> /. §31. Addition of angular momenta Let us consider a system composed of two parts whose interaction is weak. If the interaction is entirely neglected, then for each part the law of conserva- tion of angular momentum holds. The angular momentum L of the whole system can be regarded as the sum of the angular momenta Li and L2 of its parts. In the next approximation, when the weak interaction is taken into f Ordinary (polar) vectors change sign under the inversion transformation, whilst axial vectors (for instance, the vector product of two polar vectors) are unchanged by this transformation. 98 Angular Momentum §31 account, Li and L2 are not exactly conserved, but the numbers L\ and L 2 which determine their squares remain "good" quantum numbers suitable for an approximate description of the state of the system. Regarding the angular momenta in a classical manner, we can say that in this approximation Li and L2 rotate round the direction of L while remaining unchanged in magnitude. For such systems the question arises regarding the "law of addition" of angular momenta : what are the possible values of L for given values of L\ and L 2 1 The law of addition for the components of angular momentum is evident: since L z = L\ z + L 2z , it follows that M = M\ + M 2 . There is no such simple relation for the operators of the squared angular momenta, how- ever, and to derive their "law of addition" we reason as follows. If we take the quantities L-, 2 , L 2 2 , L Xz , L 2z as a complete set of physi- cal quantities,f every state will be determined by the values of the numbers L x , L 2 , M x , M 2 . For given L x and L 2 , the numbers M x and M 2 take (21^+1) and (2L 2 +1) different values respectively, so that there are altogether (2L 1 +1)(2L 2 +1) different states with the same L x and L 2 . We denote the wave functions of the states for this representation by <f> L LiMM . Instead of the above four quantities, we can take the four quantities Li 2 , L 2 2 , L 2 , L z as a complete set. Then every state is characterised by the values of the numbers L x , L 2 , L, M (we denote the corresponding wave functions by ^l^lm)- For given L x and L 2 , there must of course be (2L 1 +1)(2L 2 +1) different states as before, i.e. for given L x and L 2 the pair of numbers L and M must take (2L 1 +1)(2L 2 +1) pairs of values. These values can be determined as follows. To each value of L, there correspond 2L + 1 different possible values of M, from — L to +L. The greatest possible value of M in the states ^ (for given L x and L 2 ) is M = L x +L 2 , which is obtained when M x = L x and M 2 = L 2 . Hence the greatest possible value of M in the states iff is Lj+1,2, and this is therefore the greatest possible value of L also. Next, there are two states <f> with M = L ± +L 2 — 1, namely those where M x = L X ,M 2 — L 2 —\ and M x = L x — 1, Af 2 = L 2 . Consequently, there must also be two states ifj with this value of M ; one of them is the state with L = L x +L 2 (and M = L— 1), and the other is clearly that with L = L^+L^ — 1 (and M = L). For the value M = L x +L 2 —2 there are three different states <j>, with the following pairs of values of M lf M 2 : (L ly L 2 — 2), (L x — 1, L 2 — 1), and (L X —2,L 2 ). This means that, besides the values L = L x +L 2 , L = L x +L 2 —1, the value L = L x +L 2 — 2 can occur. The argument can be continued in this way so long as a decrease of M by 1 increases by 1 the number of states with a given M. It is easily seen that this is so until M reaches the value \L± — L 2 \. When M decreases further, the number of states no longer increases, remaining equal to 2L2+ 1 (if L 2 ^ L{). Thus \L\ — L 2 \ is the least possible value of L, and we arrive at the result f Together with such other quantities as form a complete set when combined with these four. These other quantities play no part in the subsequent discussion, and for brevity we shall ignore them entirely, and conventionally call the above four quantities a complete set. §31 Addition of angular momenta 99 that, for given L\ and L% the number L can take the values L = Li+Z^+La-l, ..., I^-L^, (31.1) that is 2L 2 +1 different values altogether (supposing that L 2 < jy. It is easy to verify that we do in fact obtain (2L 1 +1)(2L 2 +1) different values of the pair of numbers M, L. Here it is important to note that, if we ignore the 2L+\ values of M for a given L, then only one state will correspond to each of the possible values (31.1) of L. This result can be illustrated by means of what is called the vector model. If we take two vectors L l5 L 2 of lengths L x and L 2 , then the values of L are represented by the integral lengths of the vectors L which are obtained by vector addition of L x and Lg; the greatest value of L is L x -\-L 2 , which is obtained when L 2 and L 2 are parallel, and the least value is \L L —L 2 \, when Li and L 2 are antiparallel. The addition rule for angular momenta which we have obtained also makes it possible, of course, to add any number (more than two) of angular momenta by successive applications of this rule. In states with definite values of the angular momenta Z^, L 2 and of the total angular momentum L, the scalar products L x . L 2 , L . L x and L . L 2 also have definite values. These values are easily found. To calculate L x . L 2 , we write L = I^+Lg or, squaring and transposing, Replacing the operators on the right-hand side of this equation by their eigenvalues, we obtain the eigenvalue of the operator on the left-hand side: L x . L 2 = J{X(L+1)-L 1 (Z 1 + !)-£,(£,+ 1)}. (31.2) Similarly we find L . L x = ^^+l)+L 1 (L l +l)-L^L t +l)}. (31.3) Let us now determine the "addition rule for parities". As we know, the wave function T of a system consisting of two independent parts is the pro- duct of the wave functions W ± and T 2 of these parts. Hence it is clear that, if the latter are of the same parity (i.e. both change sign, or both do not change sign, when the sign of all the co-ordinates is reversed), then the wave function of the whole system is even. On the other hand, if T x and T 2 are of opposite parity, then the function Y is odd. This rule can, of course, be generalised at once to the case of a system composed of any number n of non-interacting parts. If these parts are in states with definite parities determined by the corresponding eigenvalues Ii = ± 1 of the operator /, then the parity I of the state of the whole system is given by the product /=7 1 7 2 .../ n . (31.4) 100 Angular Momentum §31 In particular, if we are concerned with a system of particles in a centrally symmetric field (the mutual interaction of the particles being supposed weak), then I t = ( — l) 1 *, where l t is the angular momentum of the *'th particle (see (30.7)), so that the parity of the state of the whole system is given by 7 = (-l)Wwtf„. (31.5) We emphasise that the exponent here contains the algebraic sum of the angular momenta l if and this is not in general the same as their "vector sum", i.e. the angular momentum L of the system. If a closed system disintegrates (under the action of internal forces), the total angular momentum and parity must be conserved. This circumstance may render it impossible for a system to disintegrate, even if this is energetic- ally possible. For instance, let us consider an atom in an even state with angular momen- tum L = 0, which is able, so far as energy considerations go, to disintegrate into a free electron and an ion in an odd state with the same angular momen- tum L = 0. It is easy to see that in fact no such disintegration can occur (it is, as we say, forbidden). For, by virtue of the law of conservation of angu- lar momentum, the free electron would also have to have zero angular momen- tum, and therefore be in an even state (I = ( — 1)° = +1); the state of the system ion + electron would then be odd, however, whereas the original state of the atom was even. CHAPTER V MOTION IN A CENTRALLY SYMMETRIC FIELD §32. Motion in a centrally symmetric field The problem of the motion of two interacting particles can be reduced in quantum mechanics to that of one particle, as can be done in classical mech- anics. The Hamiltonian of the two particles (of masses mi, mz) interacting in accordance with the law U(r) (where r is the distance between the particles) is of the form # = -— Ax-— A 2 + U(r), (32.1) 2m x 2m % where Ai arid A 2 are the Laplacian operators with respect to the co-ordinates of the particles. Instead of the radius vectors r x and r 2 of the particles, we introduce new variables R and r: r = r 2— *i> R = ( m i r i+™jr 2 )l(m x +m 2 ); (32.2) r is the vector of the distance between the particles, and R the radius vector of their centre of mass. A simple calculation gives 8 = - 0/ , An— -A + I7(r), (32.3) 2(m 1 +m 2 ) 2m v J where A R and A are the Laplacian operators with respect to the components of the vectors R and r respectively, m x +m 2 is thetotal mass of the system, and m = m 1 ni2l(m 1 +m i ) is what is called the reduced mass. Thus the Hamiltonian falls into the sum of two independent parts. Hence we can look for 0(r lf r 2 ) in the form of a product ^(R)^r(r), where the function ^(R) describes the mo- tion of the centre of mass (as a free particle of mass m^m^), and «/r(r) describes the relative motion of the particles (as a particle of mass m moving in the cen- trally symmetric field U(r)). Schrodinger's equation for the motion of a particle in a centrally sym- metric field is A0+(2m//* 2 )[£- U(r)W = 0. (32.4) Using the familiar expression for the Laplacian operator in spherical polar co-ordinates, we can write this equation in the form 1 0/ &K lr 1 d/ M\ 1 Sty! 2m (32.5) 101 102 Motion in a Centrally Symmetric Field §32 If we introduce here the operator i 2 (26.16) of the squared angular momentum we obtainf h 2 r 18/ dib\ I 2 "I £L - 7a;raVj +DW *-*- (326) The angular momentum is conserved during motion in a centrally sym- metric field. We shall consider stationary states in which l 2 and l z have definite values. In other words, we shall seek the common eigenfunctions of the operators i?, i 2 and l z . The requirement that iff is an eigenfunction of the operators i 2 and l z determines its angular dependence. We thus seek solutions of equation (32.6) in the form ^ = R(r)Y lm (9,<f>), (32.7) where the functions Y lm (6, <f>) are defined by the formulae of §28. Since l 2 Y lm = l(l+l)Y lm , we obtain for the radial function R(r) the equa- tion 1 d / dR\ 1(1+1) 2m We note that this equation does not contain the value of l z = m at all, in accordance with the (2/+l)-fold degeneracy of the levels, with which we are already familiar. Let us investigate the radial part of the wave functions. By the substitu- tion R(r) = x (r)/r (32.9) equation (32.8) is brought to the form d 2 x r2m 1(1+1)1 _ + [_ (£ _ C0 _l_l], =0 . ( 3 2 ,0) If the potential energy U(r) is everywhere finite, the wave function if/ must also be finite in all space, including the origin, and consequently so must its radial part R(r). Hence it follows that x { r ) must vanish for r = : X(0)=0. (32.11) t If we introduce the operator of the radial component p r of the linear momentum, in the form 1 2 /d 1\ prf = -ih-—(r$) = -ih( —+- U, r or \or rj the Hamiltonian can be written in the form H = (l/2m)(/ r 2 +M 2 /r 2 )+E/(r), which is the same in form as the classical Hamilton's function in spherical polar co-ordinates. §32 Motion in a centrally symmetric field 103 This condition actually holds also (see §35) for a field which becomes infinite as r -> 0. The normalisation condition for the radial function R(r) is determined by the integral J \R\ 2 r 2 dr, and therefore that for the function x(r) is determined by the integral J" |x| 2 dr. Equation (32.10) is formally identical with Schrodinger's equation for one-dimensional motion in a field of potential energy h 2 Z(Z+1) £/,(r) = £/(,)+ J_ J-, (32.12) 2m r l which is the sum of the energy U(r) and a term M(/+l)/2mr 2 = hH*l2mr\ which may be called the centrifugal energy. Thus the problem of motion in a centrally symmetric field reduces to that of one-dimensional motion in a region bounded on one side (the boundary condition for r = 0). In one-dimensional motion in a region bounded on one side, the energy levels are not degenerate ( §21). Hence we can say that, if the energy is given, the solution of equation (32.10), i.e. the radial part of the wave function, is completely determined. Bearing in mind also that the angular part of the wave function is completely determined by the values of / and m, we reach the conclusion that, for motion in a centrally symmetric field, the wave func- tion is completely determined by the values of E, I and m. In other words, the energy, the squared angular momentum and the ^-component of the angular momentum together form a complete set of physical quantities for such a motion. The reduction of the problem of motion in a centrally symmetric field to a one-dimensional problem enables us to apply the oscillation theorem (see §21). We arrange the eigenvalues of the energy (discrete spectrum) for a given / in order of increasing magnitude, and give them numbers n r , the lowest level being given the number % = 0. Then n r determines the number of nodes of the radial part of the wave function for finite values of r (excluding the point r = 0). The number nr is called the radial quantum number. The number / for motion in a centrally symmetric field is sometimes called the azimuthal quantum number, and m the magnetic quantum number. There is an accepted notation for states with various values of the angular momentum / of the particle: they are denoted by Latin letters, as follows: Z = 01234567... s p d f g h i k ... The normal state of a particle moving in a centrally symmetric field is always the s state; for, if / # 0, the angular part of the wave function in- variably has nodes, whereas the wave function of the normal state can have no nodes. We can also say that the least possible eigenvalue of the energy, 104 Motion in a Centrally Symmetric Field §33 for a given /, increases with /. This follows from the fact that the presence of an angular momentum involves the addition of the essentially positive term /* 2 /(/+l)/2mr 2 , which increases with /, to the Hamiltonian. Let us determine the form of the radial function near the origin. Here we shall suppose that lim U(r)r 2 = 0. r-*0 We seek R(r) in the form of a power series in r, retaining only the first term of the series for small r; in other words, we seek R(r) in the form R = con- stant Xr 1 . Substituting this in the equation d(r 2 d#/dr)/dr-/(/+l)i? = 0, which is obtained from (32.8) by multiplying by r 2 and taking the limit as r -> 0, we find s(s+l) =1(1+1). Hence s = I or s = — (/+1). The solution with s = — (/+1) does not satisfy the necessary conditions; it becomes infinite for r = (we recall that / ^ 0). Thus the solution with s = I remains, i.e. near the origin the wave functions of states with a given / are proportional to r l : R t £ constant xr*. (32.13) The probability of a particle's being at a distance between r and r+dr from the centre is determined by the value of r*\R\ 2 and is thus proportional to r 2(i+i) w e see t j lat - lt b ecomes 2ero at ^ or igi n tne more rapidly, the greater the value of /. §33. Free motion (spherical polar co-ordinates) The wave function of a freely moving particle t/tp = constant xc (t '/ A) P- r describes a stationary state in which the particle has a definite momentum p (and energy E = p 2 j2m). Let us now consider stationary states of a free particle in which it has a definite value, not only of the energy, but also of the absolute value and component of the angular momentum. Instead of the energy, it is convenient to introduce the wave number k=p/h = V(2mE)/h. (33.1) The wave function of a state with angular momentum / and projection §33 Free motion {spherical polar co-ordinates) 105 thereof m has the form lAwm = RM Y lm (d f <j>), (33.2) where the radial function is determined by the equation 2 r Z(/+l)n *«"+-*h'+ L* 1 """^]^ = ° ( 33 - 3 > (equation (32.8) with U(r) = 0). The wave functions ^ Wm satisfy the condi- tions of normalisation and orthogonality: / kffm+fium W = Sll'KmW-k). The orthogonality for different /, /' and m, m' is ensured by the angular func- tions. The radial functions must be normalised by the condition oo jrW^Rudr-Sik'-k). (33.4) o If we normalise the wave functions, not on the "k scale", but on the "energy scale", i.e. by the condition oo j rtR^REt dr = 8(E'-E), o then, by the general formula (5.14), we have Rbi = Ra VCdkJdE) = (l/h) V(m/k)R kl . (33.5) For / = 0, equation (33.3) can be written d 2 (r**o) dr 2 •khRn = 0; its solution finite for r = and normalised by the condition (33 4) is fcf (21.9)) v 72 sin kr . 77 r (33.6) To solve equation (33.3) with / ^ 0, we make the substitution R u = i*Xki- (33.7) For Xfci we have the equation X»"+2(/+lW/r+**«=0. If we differentiate this equation with respect to r, we obtain W" + ^," + [^-^> tt '=0. 106 Motion in a Centrally Symmetric Field §33 By the substitution xui = r Xk,i+i it becomes 2(1+2) Xfc,l+i"H Xk,l+i'+k 2 x k ,i+ 1 = 0, r which is in fact the equation satisfied by xn,i+i- Tnus the successive func- tions Xki are related by Xm+i = Xki/r, (33.8) and hence ,'1 dV Xkl ~ {;&)*»* wner e Xko = Rko is determined by formula (33.6) (this expression can, of course, be multiplied by an arbitrary constant). Thus we finally have the following expression for the radial functions in the free motion of a particle : / *s7 I 2 rl Z 1 d V sin ^ (the factor k~ l is introduced for normalisation purposes — see below — and the factor ( — l) 1 for convenience).! To obtain an asymptotic expression for the radial function (33.9) at large distances, we notice that the term which decreases least rapidly as r -» oo is obtained by differentiating sin kr I times : 12 1 d* Rki ~ (— 1) / — 77- — 7 sin kr. V 77 k l r dr l d • , / dV — smkr =ksm(kr-%7r),..., I J sinkr = # sin (Ar- A/77), t The functions R u can be expressed in terms of Bessel functions of half-integral order, in the form Ru = V(k/r)J l+1/2 (kr). (33.9a) The first few functions R kl are : \2 sinkr Rkn = l-k kr 2 rsin&r coskr" y2 rsmkr cos&r"] 7t L (kr) z k~r~J (l r/ 3 1\ 3cosArT #* 2 = -k\ ( hinkr . V77 L\(kr)* kr) (kr) 2 J §33 Free motion {spherical polar co-ordinates) 107 we have the following asymptotic expression : V2 sin(kr— Utt) — . (33.10) it r The normalisation of the functions R kl can be effected by means of their asymptotic expressions, as was explained in §21. Comparing the asymptotic formula (33.10) with the normalised function R kQ (33.6), we see that the func- tions R kl , with the coefficient used in (33.9), are in fact normalised as they should be. Near the origin (r small) we have, retaining only the term containing the lowest power of r, 1 dVsin&r /l d\*^ #2n+i r 2n /l dysin&r /ldVy #Jn+i r 2n \r drj r ~ \r drj ^ (2«+l)I /l dV^+V 2 ' = (-m — ) \rdr/(2/+l)I 2. 4... 21 = (_ l)^+i (2/+1)! = (-l)^+V1.3...(2/+l). Thus the functions i? fci near the origin have the formf in agreement with the general result (32.13). In some problems it is necessary to consider wave functions which do not satisfy the usual conditions of finiteness, but correspond to a flux of particles from the origin. The wave function which describes such a flux of particles with angular momentum / = is obtained by taking, instead of the "station- ary spherical wave" (33.6), a solution in the form of an "outgoing spherical wave", # ft0 + = Ae ikr Jr. (33.12) This function becomes infinite at the origin. Similarly, a flux of particles incident on the centre (with angular momen- tum / = 0) is described by a wave function in the form of an "ingoing spherical wave", R k0 ~ = Ae~ ik rjr. (33.13) f The symbol !! denotes the product of all integers of the same parity up to and including the num- ber in question. 108 Motion in a Centrally Symmetric Field §33 In the general case of an angular momentum / which is not zero, we obtain a solution of equation (33.3) in the formf , r 1 /l d \ l e ±ikr ^ =( _^_(._)_. (33 , +) The asymptotic expression for these functions is R kl ± « Ae ±{ < kr - ln ^Jr. (33.15) Near the origin, it has the form (2/-1)!! 22„± « A — H- 1 . (33.16) We normalise these functions so that they correspond to the emission (or absorption) of one particle per unit time. To do so, we notice that, at large distances, the spherical wave can be regarded as plane in any small interval, and the probability current density in it is i = viffif,*, where v = kh/m is the velocity of a particle. The normalisation is determined by the condition §idf =1, where the integration is carried out over a spherical surface of large radius r, i.e. j ir % do = 1, where do is an element of solid angle. If the angular functions are normalised as before, the coefficient A in the radial function must be put equal to A =1/Vv = V(mjkh). (33.17) An asymptotic expression similar to (33.10) holds, not only for the radial part of the wave function of free motion, but also for motion (with positive energy) in any field which falls off sufficiently rapidly with distance.:}: At large distances we can neglect both the field and the centrifugal energy in Schrodinger's equation, and there remains the approximate equation 1 d*(rR kl ) r dr z The general solution of this equation is R* *J IT 2 sin(Ar— ^/tt+Si) unct: rd.2), (33.18) t These functions can be expressed in terms of Hankel functions : Rm ± = ±iA V(^/2r)H?4W), (33.14a) of the first and second kinds for the signs + and — respectively. The asymptotic expansion of the functions i?*j± for large r is e ±i(kr-ln/2) p /(/ + 1) (/-l)/(/+l)(/+2) r 1 *(f+i) (f-i)*(*+i)(/+2) -j 1 l!2^r 2\(2ikrY T "j r L l!2^r 2\(2ikrf % As we shall show in §123, the field must decrease more rapidly than 1/r. §33 Free motion {spherical polar co-ordinates) 109 where S t is a constant, called the phase shift, and the common factor is chosen in accordance with the normalisation of the wave function on the "k scale".t The constant phase shift S z is determined by the boundary condition (Ria is finite as r -» 0) ; to do this, the exact Schrodinger's equation must be solved, and 8 Z cannot be calculated in a general form. The phase shifts S, are, of course, functions of both / and k, and are an important property of the eigenfunctions of the continuous spectrum. PROBLEMS Problem 1 . Determine the energy levels for the motion of a particle with angular momen- tum / = in a centrally symmetric potential well: U(r)[= — U for r < a, U(r) = for r > a. Solution. For / = the wave functions depend only on r. Inside the well, Schrodinger's equation has the form ~ TiWH*** = 0, * = - V[2m(U - |tf|)]. r dr a n The solution finite for r = is For r > a, we have the equation sinkr 1 d 2 1 " 77^)- *V - °» «=r V(2m\E\). r dr z n The solution vanishing at infinity is ifi = A'erKr/r. The condition of the continuity of the logarithmic derivative of rtf> at r = a gives k cot ka = -k = - V[(2mU /h 2 )-k 2 ], (1) or sinka = ±V(h*l2ma 2 U )ka. (2) This equation determines in implicit form the required energy levels (we must take those roots of the equation for which cot ka < 0, as follows from (1)). The first of these levels (with / = 0) is at the same time the deepest of all energy levels whatsoever, i.e. it corresponds to the normal state of the particle. If the depth U of the potential well is small enough, there are no levels of negative energy, and the particle cannot "stay" in the well. This is easily seen from equation (2), by means of the following graphical construction. The roots of an equation of the form ± sin * = ax are given by the points of intersection of the line y = oue with the curves y = ±sin x, and we must take only those points of intersection for which cot x < ; the corresponding parts of the curve y = sin x are shown in Fig. 9 by a continuous line. We see that, if a is sufficiently large (U small), there are no such points of intersection. The first such point appears when the line y = atx occupies the position Oa, i.e. for a = 2/w, and is at * = Jtt. t The term — \Itt in the argument of the sine is added so that Si = when the field is absent. Since the sign of the wave function as a whole is not significant, the phase shifts Si are determined to within nir (not 2nir). Their values may therefore always be chosen in the range between and n. 110 Motion in a Centrally Symmetric Field §33 Fig. 9 Putting a = h/V(2ma 2 U ), x = ka, we hence obtain for the minimum well depth to give a single negative level ^o.min = **h*l8ma*. (3) This quantity is the greater, the smaller the well radius a. The position of the first level Z?i at the point where it first appears is determined from ka = \tt and is E x = 0, as we should expect. As the well depth increases further, the normal level E x descends. Problem 2. Determine the order of the energy levels with various values of the angular momentum / in a very deep potential well (Uo > % % \mcF) (W. Elsasser 1933). Solution. The condition at the boundary of the well requires that -> as Uo -> oo (see §22). Writing the radial wave function within the well in the form (33.9a), we thus have the equation Jl+l/2(ka) = 0, whose roots give the position of the levels above the bottom of the well (Uo — \E\ = H 2 k 2 l2m) for various values of I. The order of the levels from the ground state is found to be 1*, lp, Id, 2s, If, 2p, \g, 2d, \h, 3s, 2/, ... . The numbers preceding the letters give the sequence of levels for each /.f Problem 3. Determine the order of appearance of levels with various I as the depth Uo of the well increases. Solution. When it first appears, each new level has energy E = 0. The corresponding wave function in the region outside the well, which vanishes asr-> oo, is Ri = constant x r~ (l +U (the solution of equation (33.3) with k = 0). From the continuity of i? ? and Ri' at the boundary of the well it follows, in particular, that the derivative (r l+x R{)' is continuous, and so we have the following condition for the wave function within the well: (ri+iRtf = for r = a. This is equivalent to the condition for the function Ri-± to vanish and, from (33.9a), we obtain the equation Ji-i/2(aV(2mU )lh) = 0; for I = the function 7?_i/ 2 must be replaced by the cosine. This gives the following order of appearance of new levels as Uo increases : 1*, lp, Id, 2s, If, 2p, \g, 2d, 3s, \h, 2f, ... . It may be noted that differences from the order of levels in a deep well occur only for compar- atively high levels. t This notation is customary for particle levels in the nucleus (see Chapter XVI). % According to (33.7) and (33.8) we have (r~ l Ri)' ~ r-'Ri +1 . Since the equation (33.3) is unaltered when / is replaced by — /— 1, we also have (r 1+1 R-i-i)' ~ r'+ii?_ ( . Finally, since the functions i?_j and Ri-i satisfy the same equation, we obtain (r l+1 Ri)' ~ r l+1 Ri_i, the formula used in the text. §34 Resolution of a plane wave 111 Problem 4. Determine the energy levels of a three-dimensional oscillator (a particle in a field U — $fi(o*r 2 ), their degrees of degeneracy, and the possible values of the orbital angular momentum in the corresponding stationary states. Solution. Schrodinger's equation for a particle in a field U = $na> 2 (x*+y t +z*) allows separation of the variables, leading to three equations like that of a linear oscillator. The energy levels are therefore E n = foofa+wj+ns+f) = ha>{n+f). The degree of degeneracy of the nth level is equal to the number of ways in which n can be divided into the sum of three positive integral (or zero) numbers ;f this is K»+i)(»+2). The wave functions of the stationary states are «£ WlW ,n 3 = constant x e-^V2H n ^)H n J < pLy)H n3 {a.z), (1) where a = \/(ma>lfi) and m is the mass of the particle. When the sign of the co-ordinate is changed, the polynomial H n is multiplied by ( — 1)". The parity of the function (1) is therefore ( — l)»i+»s+»8 = ( — 1)". Taking linear combinations of these functions with a given sum ni + m + nz = n, we can form the functions 4>nim = constant x e-* 2 r 2 l*r n ®i m {d)e ±im <l>, (2) where m = 0,1, ..., / and I takes the values 0, 2, ..., n for even n and 1, 3, ..., « for odd n. This is evident from a comparison of the parities ( — 1)" of the functions (1) and ( — I) 1 of the functions (2), which must be the same. This determines the possible values of the orbital angular momentum corresponding to the energy levels considered. The order of levels of the three-dimensional oscillator is, therefore, with the same notation as in Problems 2 and 3, (Is), (l/>), (Id, 2.), (1/, 2p), (\g, 2d, 3s), ... , where the parentheses enclose sets of degenerate states. §34. Resolution of a plane wave Let us consider a free particle moving with a given momentum p = kk in the positive direction of the ^-axis. The wave function of such a particle is of the form iff = constant xe ikz . Let us expand this function in terms of the wave functions rji klm oi free motion with various angular momenta. Since, in the state considered, the energy has the definite value k 2 h 2 ]2m, it is clear that only functions with this k will appear in the required expansion. Moreover, since the function e ikz has axial symmetry about the ^-axis, its expansion can contain only functions independent of the angle <f>, i.e. functions with m = 0. Thus we must have t In other words, this is the number of ways in which n similar balls can be distributed among three urns. 112 Motion in a Centrally Symmetric Field §34 where the ai are constants. Substituting the expressions (27.3), (28.7) and (33.9) for the functions <I>, 0, R, we obtain v^ /r\ l /l dVsinAr e , te= gc,p,(-*)(-)(--)_ <»-»»•>. where the C, are other constants. These constants are conveniently deter- mined by comparing the coefficients of (r cos 6) n in the expansions of the two sides of the equation in powers of r. On the right-hand side of the equation this term occurs only in the nth summand ; for / > n, the expansion of the radial function begins at a higher power of r, while for / < n the polynomial P i (cos 6) contains only lower powers of cos 0. The term in cos * 6 in P, (cos 6) has the coefficient (2/)!/2*(/!) 2 (see formula (c.2)). Using also formula (33.11), we find the desired term of the expansion of the right-hand side of the equation to be 2'(Z!) 2 (2/+1)H On the left-hand side of the equation the corresponding term (in the expansion of e ikr cos d ) is (ikr cos ey III Equating these two quantities, we find C, = (— *y(2/+l). Thus we finally obtain the required expansion: <- -g^^+DIM^Q (--) — . (34.1) At large distances this relation takes the asymptotic form e ikz ~ _ y ;/( 2 /+ l)P,(cos 6) sm(kr-ilir). (34.2) kr r—' We normalise the wave function e ik * to give a probability current density of unity, i.e. so that it corresponds to a flux of particles (parallel to the #-axis) with one particle passing through unit area in unit time. It is easy to see that this function is xfj = v -i/2 e tkz = v(m/&#)e** z> (34.3) where v is the velocity of the particles. Multiplying both sides of equation (34.1) by y/{mjkh) and introducing on the right-hand side the normalised functions ^ ffl ± = R kl ±(r)Y lm (d,<f>), we obtain 1 <» 1 —em = 2 VK2/+l)]-(^ ro +-^ -). §35 "Fall" of a particle to the centre 1 13 The squared modulus of the coefficient of ip kW ~ (or <p kt0 +) in this expansion determines, according to the usual rules, the probability that a particle in a current converging to (or diverging from) the centre has an angular momen- tum / (about the origin). Since the wave function «rV** corresponds to a current of particles of unit density, this "probability" has the dimensions of length squared; it can be conveniently interpreted as the magnitude of the "cross-section" (in the *y-plane) on which the particle must fall if its angular momentum is /. Denoting this quantity by a h we have ex, = ir(2l+ 1W = *W2ir)«(2/+ 1), (34.4) where A is the de Broglie wavelength of the particle. For large values of /, the sum of the cross-sections over a range A/ of I (such that 1 <^ A/ < /) is 2 it ih 2 <t, £ — 2/AZ = 2tt — A/. Al R P On substituting the classical expression for the angular momentum, hi = pp (where p is what is called the impact parameter), this expression becomes 27TpAp t in agreement with the classical result. This is no accident; we shall see below that, for large values of /, the motion is quasi-classical (see Chapter VII). §35. "Fall" of a particle to the centre To reveal certain properties of quantum-mechanical motion it is useful to examine a case which, it is true, has no direct physical meaning: the motion of a particle in a field where the potential energy becomes infinite at some point (the origin) according to the law U(r) « -£/r 2 , > 0; the form of the field at large distances from the origin is here immaterial. We have seen in §18 that this is a case intermediate between those where there are ordinary stationary states and those where a "fall" of the particle to the origin takes place. Near the origin, Schrodinger's equation in the present case is R"+2R'lr+yRlr* = 0, (35.1) where R(r) is the radial part of the wave function, and we have introduced the constant Y «2«0/*»-Z(Z+l) (35.2) and have omitted all terms of lower orders in 1/r; the value of the energy E is supposed finite, and so the corresponding term in the equation is omitted also. 114 Motion in a Centrally Symmetric Field §35 Let us seek R in the form R ~r s ; we then obtain for s the quadratic equation *(*+l)+y =0, which has the two roots *i = -B- V(i-y), *, = -*-V(i-y). (35.3) For further investigations it is convenient to proceed as follows. We draw a small region of radius r round the origin, and replace the function — yjr 2 in this region by the constant — y/r 2 . After determining the wave functions in this "cut off" field, we then examine the result of passing to the limit 'o->0. Let us first suppose that y < \. Then s x and s 2 are real negative quantities, and s x > s 2 . For r > r Qy the general solution of Schrodinger's equation has the form (always restricting ourselves to small r), R = Ar\+Br\ (35.4) A and B being constants. For r < r , the solution of the equation R"+2R'/r+yRlr * = which is finite at the origin has the form sin&r R = C , k = Vylr . (35.5) r For r = r , the function R and its derivative R' must be continuous. It is convenient to write one of the conditions as a condition of continuity of the logarithmic derivative of rR. This gives the equation ^i+lK Sl +£(*2+l)V' t 7 = k cot kr . i4r *»+ 1 +-Br ».+ 1 or = Vy cot \/y. Ar **+Br *> On solving for the ratio B/A, this equation gives an expression of the form B/A = constant xr s ^-^. (35.6) Passing now to the limit r -> 0, we find that BjA -*• (recalling that h > h)- Thus, of the two solutions of Schrodinger's equation (35.1) which diverge at the origin, we must choose that which becomes infinite less rapidly: R = A/rW. (35.7) § 3 *> "Fall" of a particle to the centre 115 Next, let y > £. Then s x and s 2 are complex: *i = -i+*V(y-£), * 2 = V- Repeating the above analysis, we again arrive at equation (35.6), which, on substituting the values of ^ and s 2 , gives B/A = constant xr^^H. (35.8) On passing to the limit r -» 0, this expression does not tend to any definite limit, so that a direct passage to the limit is not possible. Using (35.8), the general form of the real solution can be written R = constant xrW cos( V(y-£) log (r/r )+ constant). (35.9) This function has a number of zeros which increases without limit as r decreases. Since, on the one hand, the expression (35.9) is valid for the wave function (when r is sufficiently small) with any finite value of the energy E of the particle, and, on the other hand, the wave function of the normal state can have no zeros, we can infer that the "normal state" of a particle in the field considered corresponds to the energy E = - oo. In every state of a discrete spectrum, however, the particle is mainly in a region of space where E > U. Hence, for E -> - oo, the particle is in an infinitely small region round the origin, i.e. the particle "falls" to the centre. The "critical" field U cr for which the "fall" of a particle to the centre becomes possible corresponds to the value y = J. The smallest value of the coefficient of -1/r 2 is obtained for / = 0, i.e. Ucr = -h 2 /8mr 2 . (35.10) It is seen from formula (35.3) (for s x ) that the permissible solution of Schro- dinger's equation (near the point where U ~ 1/r 2 ) diverges, as r -> 0, not more rapidly than 1/yV. If the field becomes infinite, as r -> 0, more slowly than 1/r 2 , we can neglect U{r), in Schrodinger's equation near the origin, in comparison with the other terms, and we obtain the same solutions as for free motion, i.e. if, ~ r l (see §33). Finally, if the field becomes infinite more rapidly than 1/r 2 (as -1/r 8 with s > 2), the wave function near the origin is proportional to r**-i (see §49, Problem). In all these cases the product rip tends to zero at r = 0. Next, let us investigate the properties of the solutions of Schrodinger's equation in a field which diminishes at large distances according to the law U « -£/r 2 , and has any form at small distances. We first suppose that y <h It is easy to see that in this case only a finite number of negative energy levels can exist.f For with energy E = Schrodinger's equation at large distances has the form (35.1), with the general solution (35.4). The function (35.4), however, has no zeros (for r # 0); hence all zeros of the required radial wave function lie at finite distances from the origin, and their t It is assumed that for small r the field is such that the particle does not "fall". 5 116 Motion in a Centrally Symmetric Field §36 number is always finite. In other words, the ordinal number of the level E = which terminates the discrete spectrum is finite. If y > I, on the other hand, the discrete spectrum contains an infinite number of negative energy levels. For the wave function of the state with E = has, at large distances, the form (35.9), with an infinite number of zeros, so that its ordinal number is always infinite. Finally, let the field be U = -jS/r 2 in all space. Then, for y > £, the particle "falls", but if y < £ there are no negative energy levels. For the wave function of the state with E = is of the form (35.7) in all space ; it has no zeros at finite distances, i.e. it corresponds to the lowest energy level (for the given /). §36. Motion in a Coulomb field (spherical polar co-ordinates) A very important case of motion in a centrally symmetric field is that of motion in a Coulomb field U = ±«/r (where a is a positive constant). We shall first consider a Coulomb attraction, and shall therefore write U = —<xjr. It is evident from general considera- tions that the spectrum of negative eigenvalues of the energy will be discrete (with an infinite number of levels), while that of the positive eigenvalues will be continuous. Equation (32.8) for the radial functions has the form d 2 R 2dR 1(1+1) nt 2m/ — —- 1 — 2m/ a\ .#+_(£+-)# =0. (36.1) H 2 \ rj dr 2 r dr If we are concerned with the relative motion of two attracting particles, m must be taken as the reduced mass. In calculations connected with the Coulomb field it is convenient to use, instead of the ordinary units, special units for the measurement of all quanti- ties, which we shall call Coulomb units. As the units of measurement of mass, length and time, we take respectively m, h 2 /mx, ft? I ma. 2 . All the remaining units are derived from these ; thus the unit of energy is mx 2 /ft 2 . From now on, in this section and the following one, we shall always (unless explicitly stated otherwise) use these units.f t Ifm = 911xl0 _28 gis the mass of the electron, anda = e 2 (where e is the charge on the electron), the Coulomb units are the same as what are called atomic units. The atomic unit of length is ft 2 \me 2 = 0-529 x lO" 8 cm (what is called the Bohr radius). The atomic unit of energy is me^lh 2 = 4-36 XIO" 11 erg = 27-21 electron-volts. The atomic unit of charge is e = 4-80 X 10 -10 esu. We formally obtain the formulae in atomic units by putting e = m = H = 1 . For a = Ze 2 the Coulomb and atomic units are not the same. §36 Motion in a Coulomb field {spherical polar co-ordinates) 117 Let us rewrite equation (36.1) in the new units: d 2 R 2 dR 7(7+1) / 1\ Instead of the parameter E and the variable r, we introduce the new quantities n = l/V(-2£), P = 2r/n. (36.3) For negative E (which we shall first consider), n is a real positive number. The equation (36.2), on making the substitutions (36.3), becomes 2 r n 7(7+1) n #"+-#'+ -£+— - L \R=o (36.4) P L p p 2 J (the primes denote differentiation with respect to p). For small p, the solution which satisfies the necessary conditions of finite- ness is proportional to p l (see (32.13)). To calculate the asymptotic be- haviour of R for large p, we omit from (36.4) the terms in 1/p and 1/p 2 and obtain the equation R" = IR, whence R = e ±iP . The solution in which we are interested, which vanishes at infinity, consequently behaves as e~ ip for large p. It is therefore natural to make the substitution R = P l e- P l*w{p\ (36.5) when equation (36.4) becomes pw"+(2l+2-p)w'+(n-l-l)w = 0. (36.6) The solution of this equation must diverge at infinity not more rapidly than every finite power of p, while for p = it must be finite. The solution which satisfies the latter condition is the confluent hypergeometric function to = F(-n+l+l, 27+2, p) (36.7) (see §d of the Mathematical Appendices).f A solution which satisfies the condition at infinity is obtained only for negative integral (or zero) values of — n+7+1, when the function (36.7) reduces to a polynomial of degree n—l—1. Otherwise it diverges at infinity as eP (see (d.14)). Thus we reach the conclusion that the number n must be a positive integer, and for a given 7 we must have n>l+l. (36.8) Recalling the definition (36.3) of the parameter n, we find E = -l/2n\ n = l,2,.... (36.9) t The second solution of equation (36.6) diverges as p -2 ' -1 as p -> 0. 118 Motion in a Centrally Symmetric Field §36 This solves the problem of determining the energy levels of the discrete spectrum in a Coulomb field. We see that there are an infinite number of levels between the normal level Ei = — \ and zero. The distances between successive levels diminish as n increases; the levels become more crowded as we approach the value E = 0, where the discrete spectrum closes up into the continuous spectrum. In ordinary units, formula (36.9) isf E = -ma?l2h*n 2 . (36.10) The integer n is called the principal quantum number. The radial quantum number defined in §32 is n r = n—l—l. For a given value of the principal quantum number, / can take the values Z = 0,l,...,w-1, (36.11) i.e. n different values in all. Only n appears in the expression (36.9) for the energy. Hence all states with different / but the same n have the same energy. Thus each eigenvalue is degenerate, not only with respect to the magnetic quantum number m (as in any motion in a centrally symmetric field) but also with respect to the number /. This latter degeneracy (called accidental) is a specific property of the Coulomb field. To each value of / there correspond, as we know, 2/+1 different values of m. Hence the degree of degeneracy of the nth. energy level is S X (2/+1)=« 2 . (36.12) 1=0 The wave functions of the stationary states are determined by formulae (36.5) and (36.7). The confluent hypergeometric functions with both parameters integral are the same, apart from a factor, as what are called the generalised Laguerre polynomials (see §d of the Mathematical Appendices). Hence _ 7 /or 2I+l R n i = constant xp l e-<>' 2 L n +i (p). The radial functions must be normalised by the condition j R n i 2 r 2 dr = 1. t Formula (36.10) was first derived by N. Bohr in 1913, before the discovery of quantum mechanics. In quantum mechanics it was derived by W. Pauli in 1926 using the matrix method, and a few months later by E. Schrodinger using the wave equation. §36 Motion in a Coulomb field {spherical polar co-ordinates) 119 Their final form isf 2 / («+/) / (n+l)\ V (n-/-1)! ( 2 ^- r/ ^(-«+/+l, 2/+ 2, 2r/n) ; «'+ 2 (2/+l)! V(n-/-l)I (36.13) the normalisation integral is calculated by (f.6)4 Near the origin, R ni has the form At large distances, R nl « (-l)n-l-l r n-l e -r/n ,35 15) The wave function R 10 of the normal state decreases exponentially at distances of the order r ~ 1, i.e. r ~ h^jnix in ordinary units. The mean values of the various powers of r are calculated from the formula 00 r k = J* ^+2^2 dr< The general formula for r k can be obtained by means of formula (f.7). Here we shall give the first few values of r* (for positive and negative k) : r =i[3n»-/(/+l)], r* =4»*[5»*+l-3/(/+l)] f ^ =r =l/»2, r "=i = l/„3(/ + J). (36.16) t We give the first few functions R nl explicitly: #20=(1/V2>-^(1-^), R 21 =(l/2V6)e-r/2 r , R 30 = (2/3 v3)e-'/3A--r+-A fl 31 = (8/27V6)e-^/3/l— X R 32 = (4/81 V30)e-r/3 r 2. X The normalisation integral can also be calculated by substituting the expression (d.13) for the Laguerre polynomials and integrating by parts (similarly to the calculation of the integral (c.ll) for the Legendre polynomials). 120 Motion in a Centrally Symmetric Field §36 The spectrum of positive eigenvalues of the energy is continuous and extends from zero to infinity. Each of these eigenvalues is infinitely degener- ate ; to each value of E there corresponds an infinite number of states, with / taking all integral values from to oo (and with all possible values of m for the given /). The number n and the variable p, defined by the formulae (36.3), are now purely imaginary : n = _;/v(2E) = -ijk, P = 2ikr (36.17) (we have introduced the wave number k = y/(2E) in place of the energy). The radial eigenfunctions of the continuous spectrum are of the form R» = —^—(2krye-i k 'F(ilk+l+l, 21+2, 2ikr), (36.18) where the C k are normalisation factors. They can be represented as a complex integral (see §d) : 1 r / 2ikr\- i l k - l -' i , R M = C^krfe^— i> et(l —J t*~ 2 df, (36.19) which is taken along the contourf shown in Fig. 10. The substitution ,, i-zikr Fig. 10 | = 2ikr(t + £) converts this integral to the more symmetrical form (-2kr)~ 1 - 1 r R kl = c k - — &> <* l * r *(t+l) i / k -*- 1 (t-t)-*' k -i- 1 dt; (36.20) the path of integration passes in the positive direction round the points t = ± \. It is seen at once from this representation that the functions R kl are real. The asymptotic expansion (d.14) of the confluent hypergeometric function enables us to obtain immediately a similar expansion for the wave functions R kt . The two terms in (d.14) give two complex conjugate expressions in the | Instead of this contour we could use any closed loop passing round the singular points £ = and £ = 2ikr in the positive direction. For integral /, the function V(g) = ^ Jn - l ^—2ikr) n ^ (see §d) returns to its initial value on passing round such a contour. §36 Motion in a Coulomb field {spherical polar co-ordinates) 121 function R hh and as a result we obtain e -"/2k t e -Hkr-irQ+i)/ 2+(l Ik) log 2kr] Rh = C k — — re G(/+l+i/*, i/*-Z, -2ftr) &r ( r(Z+l— */&) (36.21) If we normalise the wave functions on the "k scale" (i.e. by the condition (33.4)), the normalisation coefficient is C h = V(2/7r)fo"/ 2 *|r(Z+l-*7A)|. (36.22) For the asymptotic expression for R kl when r is large (the first term of the expansion (36.21)) is then of the form /2 1 1 R ki ~ / - -sin(Ar+- log 2kr— |Ztt+Sj), V irr k (36.23) 8,=argr(/+l-*/*), in agreement with the general form (33.18) of the normalised wave functions of the continuous spectrum in a centrally symmetric field. The expression (36.23) differs from (33.18) by the presence of a logarithmic term in the argu- ment of the sine ; however, since log r increases only slowly compared with r itself, the presence of this term is immaterial in calculating a normalisation integral which diverges at infinity. The modulus of the gamma function which appears in the expression (36.22) for the normalisation factor can be expressed in terms of elementary functions. Using the familiar properties of gamma functions: we have and also r(*+l) =zT(z), r(s)r(l-*) =7r/sin7rs, r(/+i+f/*) = (/+»/*) ... (i+»/*)(i/ft)r(i/*), ry+i-i/k) = (i-ilk) ... (i-i/k)r(i-i/k), |r(/+i-t/ft)| = [r(/+i-^)r(/+i+^)] 1/2 -J\i\J(?+i)***l Thus 2Vk V(l-g-2"/*) for / = the product is replaced by unity tlJi***)'' (3624) 122 Motion in a Centrally Symmetric Field §36 The radial functions R El normalised on the "energy scale" are obtained from the functions R kl by dividing by \/k: R El = k~ i R Jel (see (33.5)). By passing to the limit as E -» (i.e. k -> 0), we can correctly obtain from R m the normalised radial function R ol for the particular case of zero energy, f The limit of the series F(ijk+l+l, 2/ +2, likr) as k -> is r% try \« 1 — + — ... = (2/+l)!(2r)-*-i/2/ 2l+1 ( V[8r]), (2/+ 2)1! (2/+2)(2/+3)2! where J21+1 is the Bessel function of order 21+ 1. The coefficient C& (36.24) behaves as 2k~ l+1/2 as k -> 0. We hence easily obtain R Ql = V(2/r)J 2l+1 (V[8r]). (36.25) The asymptotic form of this function for large r isj R Ql « (2/7rV)i/4 S in(v[8r]-/7r-j7r). (36.26) In a repulsive Coulomb field (U = a/r) there is only a continuous spectrum of positive eigenvalues of the energy. Schrodinger's equation in this field can be formally obtained from the equation for an attractive field by changing the sign of r. Hence the wave functions of the stationary states are found immediately from (36.18) by the same alteration. The normalisation co- efficient is again determined from the asymptotic expression, and as a result we obtain R H = —j^—(2kr)¥^F(ilk+l+l, 21+2, -2ikr) t 2 -ke-"l* k \T(l+ l+i/k)\ = v(* 2 */*-i) \ J v V + *v (36 ' 27) The asymptotic expression for this function for large r is Rm ~ /--sinf &-— -log2&r— J/w+8,J, B t =ar g r(/+l +*y&). (36.28) PROBLEMS Problem 1. Determine the probability distribution of various values of the momentum in the ground state of the hydrogen atom. t It is found that in fact the function RBi normalised on the energy scale remains finite as E -> 0, while Rki -> as k -*■ 0. % It may be noted that this function corresponds to the quasi-classical approximation (§49) applied to motion in the region (/+i) 2 <^ r <^ k~ 2 . §36 Motion in a Coulomb field {spherical polar co-ordinates) 123 Solution.! The wave function of the ground state is «A = R10Y00 = (l/\Z"-)«~ r . The wave function of this state in the p representation is then given by the integral a(p) = (2tt)-3/2 f 0(r)<r*p.r dV (see (15.10)). The integral is calculated by changing to spherical polar co-ordinates with the polar axis along p; the result is$ «(p) = 77 (1+^ 2 ) 2 and the probability density in p-space is |a(p)| 2 . Problem 2. Determine the mean potential of the field created by the nucleus and the electron in the ground state of the hydrogen atom. Solution. The mean potential <f> e created by an "electron cloud" at an arbitrary point r is most simply found as the spherically symmetric solution of Poisson's equation with charge density p = — \ifi\ 2 : 1 d 2 —ri r h) = 4e ~ 2r - r dr 1 Integrating this equation, and choosing the constants so that <f> e (0) is finite and <£ e (oo) = (), and adding the potential of the field of the nucleus, we obtain t = -+M r ) = (-+iV 2r - For r^lwe have <f> ~ 1/r (the field of the nucleus), and for r > 1 the potential 4> ^ e~ 2r (the nucleus is screened by the electron). Problem 3. Determine the energy levels of a particle moving in a centrally symmetric field with potential energy U = A/r 2 —B/r (Fig. 11). Solution. The spectrum of positive energy levels is continuous, while that of negative levels is discrete; we shall consider the latter. Schrodinger's equation for the radial func- tion is d 2 i? 2dR 2m/ h* 1 A B\ + +— ( E /(/+ 1) +— )R = 0. (1) dr 2 r dr h*\ 2rn V r* r J l ' We introduce the new variable p =2V(-2mE)r/h, and the notation 2mAJh a +l(l+l) =s(s+l), (2) BV(m/-2E)Jh =n. (3) f In Problems 1 and 2, atomic units are used. % The wave functions in the p representation for excited states in a Coulomb field are given by H. A. Bethe and E. E. Salpeter, Handbuch der Physik 35, 88, Springer, Berlin 1957. These functions were analysed by V. A. Fok and applied by him to calculate a number of complicated sums (Izvestiya Akademii Nauk SSSR, Seriya fizicheskaya No. 2, 169, 1935). 124 Motion in a Centrally Symmetric Field u(r) §36 Fig. 11 Then equation (1) takes the form 2 P 2 /In s(s+ 1)\ #"+-*'+( --+— ^— - )R = 0, P \ 4 p p 2 / which is formally identical with (36.4). Hence we can at once conclude that the solution satisfying the necessary conditions is R = p» e -Pl*F(-n+s+l, 2s+2, p), where n—s — 1 = p must be a positive integer (or zero), and s must be taken as the positive root of equation (2). From the definition (3) we consequently obtain the energy levels ■E n = 2B 2 m h* -RM-1+ V{(2/+l) 2 4-8m,4//*2}]-2. Problem 4. The same as Problem 3, but with U = A/^+Br 2 (Fig. 12). Solution. There is only a discrete spectrum. Schrodinger's equation is d 2 R 2dR 2m r k 2 l(l+l) A 1 1 — I E Br 2 dr 2 r dr h 2 L 2mr 2 r 2 ]*- 0. Introducing the variable | = V(2mB)r i /H Fig. 12 §37 Motion in a Coulomb field {parabolic co-ordinates) 125 and the notation /(/ + 1) + 2m,4//* 2 = 2s(2s+l), V(2m/B)Elh = 4(n+s)+3, we obtain the equation #r+-ir+[n+H-|-tf-*(*+i)/fl* = o. The solution required behaves asymptotically as e~$£ when £ -> co, while for small f it is proportional to £*, where s must be taken as the positive quantity s = i[-l+ V{(2/+l) 2 +8m^ 2 }]. Hence we seek a solution in the form R = e-UH'zo, obtaining for w the equation £w"+(2s+ — nw'+nw = 0, whence w = F(-n,2s+-^y where n must be a non-negative integer. We consequently find as the energy levels the infinite set of equidistant values E n =^V(5/2m)[4n+2+V{(2/+l) 2 +8m^ 2 }],n =0,1,2,.... §37. Motion in a Coulomb field (parabolic co-ordinates) The separation of the variables in Schrodinger's equation written in spherical polar co-ordinates is always possible for motion in any centrally symmetric field. In the case of a Coulomb field, the separation of the variables is also possible in what are called parabolic co-ordinates. The solution of the problem of motion in a Coulomb field in terms of parabolic co-ordinates is useful in investigating a number of problems where a certain direction in space is distinctive; for example, for an atom in an external electric field (see §77). The parabolic co-ordinates £, 77, (f> are defined by the formulae x = V(£})W, y = V(£})sin<£, * =l(£—q), r = v(* 2 +:y 2 +2 2 ) = &£+!?), or conversely |=r+*, r, =r-z, <j>=tBir 1 {ylxy, (37.2) f and rj take values from to 00, and <j> from to 2tt. The surfaces £ = 126 Motion in a Centrally Symmetric Field §37 constant and rj = constant are paraboloids of revolution about the z-sotis, with focus at the origin. This system of co-ordinates is orthogonal. The element of length is given by the expression (dZ)* = ^(d^+^Cd^+^d^, (37.3) and the element of volume is dV = £(£+77)d£Vty. (37.4) From (37.3) we have the Laplacian operator 4 rd / d\ d / d\~] 1 8* A= sAii(^K(%)]w (375) Schrodinger's equation for a particle in an attractive Coulomb field with U = -1/r = -2/(| + i7) is 4 rd/ddt\ d / 8ifr\-i 1 2Ni / 2 \ irfii(y + 4wJ + i^ + r w - °- (376) Let us seek the eigenfunctions «/r in the form =mm-ny im \ (37.7) where m is the magnetic quantum number. Substituting this expression in equation (37.6) multiplied by £(£+*?)» and separating the variables £ and rj, we obtain for/i and/2 the equations d/ d/i\ ^-(^)+[^-Jm^+)8 2 ]/ 2 =0, (37.8) diy \ d?7 where the separation parameters j8i, /?2 are related by ft+Ai = 1. (37.9) Let us consider the discrete energy spectrum (E < 0). We introduce in place of E, £, 77 the quantities « = 1IV(-2E), Px = €V(-2E) = £ln, p 2 = ,/», (37.10) whereupon we obtain the equation for /1 "P7+-— + -!+-( — — +«i )-— /1 = 0, (37.11) d/>i 2 Pi d Pl L Pi \ 2 / 4 Pl 2 J §37 Motion in a Coulomb field {parabolic co-ordinates) 127 and a similar equation for/2, with the notation "1 = -MM + l)+"ft, n 2 = -£(|m| + l)+njS 2 . (37.12) Similarly to the calculation for equation (36.4), we find thatyj behaves as e'tP 1 for large p x and as p ± iw for small p x . Accordingly, we seek a solution of equation (37.11) in the form A(Pi) = r^w^^, and similarly for/2, obtaining for zoi the equation ■Pi«'i"+(M + 1 — Pi) w i+ n i w i =°- This is again the equation for a confluent hypergeometric function. The solution satisfying the conditions of finiteness is wi =F(—n 1 , |w| + l,p x ), where n\ must be a non-negative integer. Thus each stationary state of the discrete spectrum is determined in para- bolic co-ordinates by three integers : the parabolic quantum numbers «i and «2, and the magnetic quantum number m. For n, the principal quantum number, we have from (37.9) and (37.12) n =w 1 +« 2 +|m| + l. (37.13) For the energy levels, of course, we obtain our previous result (36.9). For given «, the number \m\ can take n different values from to n — \. For fixed n and \m\ the number n x takes n—\m\ values, from to n— \m\ — 1. Taking into account also that for given \m\ we can choose the functions with m = ±N|, we find that for a given n there are altogether n-l 2 2 («-ot)+(k-0) = w 2 m—l different states, in agreement with the result obtained in §36. The wave functions ifin^m of the discrete spectrum must be normalised by the condition 00 00 2W j I^J 2 dV = J JJJ l-An^l 2 ^) d^d, = 1. (37.14) 000 The normalised functions are V2 / £\ /*?\ e im * *.%- - *A»(jJA-kh^ (37 ' 15) where 1 /(/>+|m|)! fpm(p) = — / ^ (-A M + 1. /»)^V""». (37.16) 128 Motion in a Centrally Symmetric Field §37 The wave functions in parabolic co-ordinates, unlike those in spherical polar co-ordinates, are not symmetrical about the plane z = 0. For n± > w 2 the probability of finding the particle in the direction z > is greater than that for z < 0, and vice versa for n\ < n%. To the continuous spectrum (E > 0) there corresponds a continuous spec- trum of real values of the parameters ft, ft in equations (37.8) (connected as before, of course, by the relation (37.9)). We shall not pause to write out here the corresponding wave functions, since it is not usually necessary to employ them. Equations (37.8), regarded as equations for the "eigenvalues" of the quantities ft, ft, have also (for E > 0) a spectrum of complex values of ft and ft. The corresponding wave functions are written out in §133, where we shall use them to solve a problem of scattering in a Coulomb field. In classical motion of a particle in a Coulomb field there is a conservation law peculiar to this type of field :f A = pxl — x\r = constant. (37.17) In quantum mechanics the three components of this vector cannot simul- taneously have definite values, since the operators^, A y , A z do not commute. Any one of these operators, A z say, commutes (like the ^-component of any vector; see (29.3)) with l z , but does not commute with the conserved square of the angular momentum, I 2 . The existence of another conserved quantity which does not commute with the others leads (see §10) to an additional degeneracy of the levels, and this is the accidental degeneracy peculiar to a Coulomb field. The description of motion in a Coulomb field by means of the wave functions ijt nlm in spherical polar co-ordinates corresponds to states in which not only the energy but also the square of the angular momentum and its sr-component have definite values. The wave functions ifj ni n,n in parabolic co-ordinates, on the other hand, describe stationary states in which l z and A z have definite values. It can be shown that the value of A z is given in terms of the quantum numbers n\, n%, n by A z = (nz-n{)\n. (37.18) t See Mechanics, §15. CHAPTER VI PERTURBATION THEORY §38. Perturbations independent of time The exact solution of Schrodinger's equation can be found only in a com- paratively small number of the simplest cases. The majority of problems in quantum mechanics lead to equations which are too complex to be solved exactly. Often, however, quantities of different orders of magnitude appear in the conditions of the problem ; among them there may be small quantities such that, when they are neglected, the problem is so much simplified that its exact solution becomes possible. In such cases, the first step in solving the physical problem concerned is to solve exactly the simplified problem, and the second step is to calculate approximately the errors due to the small terms that have been neglected in the simplified problem. There is a general method of calculating these errors; it is called perturbation theory. Let us suppose that the Hamiltonian of a given physical system is of the form where V is a small correction {ox perturbation) to the unperturbed operator i? . In §§38, 39 we shall consider perturbations V which do not depend explicitly on time (the same is assumed regarding i? also). The conditions which are necessary for it to be permissible to regard the operator V as "small" com- pared with the operator fi will be derived below. The problem of perturbation theory for a discrete spectrum can be formu- lated as follows. It is assumed that the eigenfunctions «/r n (0) and eigenvalues E n w of the discrete spectrum of the unperturbed operator ff are known, i.e. the exact solutions of the equation #«# > = &°ty<» (38.1) are known. It is desired to find approximate solutions of the equation ify= (#0+^)0=^, (38.2) i.e. approximate expressions for the eigenfunctions ift n and eigenvalues E n of the perturbed operator i?. In this section we shall assume that no eigenvalue of the operator i? is degenerate. Moreover, to simplify our results, we shall suppose that there is only a discrete spectrum of eigenvalues ; all the formulae can be at once generalised to the case where there is a continuous spectrum. The calculations are conveniently performed in matrix form throughout 129 130 Perturbation Theory §38 To do this, we expand the required function $ in terms of the functions &* (0) : t = Zc m i/, m «». (38.3) Substituting this expansion in (38.2) we obtain multiplying both sides of this equation by i/; k W* and integrating, we find {E-E^)c k = S V km c m . (38.4) Here we have introduced the matrix V km of the perturbation operator V, defined with respect to the unperturbed functions <^ m (0) : V km =j^*^ m (o) dq , (38>5) We shall seek the values of the coefficients c m and the energy E in the form of series E = #»+jEa> +jE » + ... , Cm = cj»+c m u+cj»+ where the quantities E™ and cJU are of the same order of smallness as the perturbation P, the quantities £< 2 > and c m < 2 > are of the second order of small- ness (if V is of the first order), and so on. Let us determine the corrections to the nth. eigenvalue and eigenfunction, putting accordingly * B <°> = 1, c m <°> = for m # n. To find the first approxi- mation, we substitute in equation (38.4) E = E n ^+E n ( x \ c k = c k W+cJ*\ and retain only terms of the first order. The equation with k = n gives E n {1) = V nn =j +®*fy® dq. (38.6) Thus the first-order correction to the eigenvalue £M°> is equal to the mean value of the perturbation in the state tp n (0) . The equation (38.4) with k ^ n gives '* (1) = V^KE^-E^) for k^n, while c n <u remains arbitrary; it must be chosen so that the function tp n = *Pn 0) + l f'n (1) is normalised up to and including terms of the first order. For this we can put c TC (1 > = 0. For the functions m ^n ■L'm (the prime means that the term with m = n is omitted from the sum) are orthogonal to n <«>, and hence the integral of l^^+^l 2 differs from unity only by a quantity of the second order of smallness. §38 Perturbations independent of time 131 Formula (38.7) determines the correction to the wave functions in the first approximation. Incidentally, we see from this formula the condition for the applicability of the above method of perturbation theory. This condition is that the inequality \V mn \ <\E®-EJ>\ (38.8) must hold, i.e. the matrix elements of the operator ^must be small compared with the corresponding differences between the unperturbed energy levels. Next, let us determine the correction to the eigenvalue £ n <°> in the second approximation. To do this, we substitute in (38.4) E = E n <°)+E n M+E (2) c k = £fc (0) +£ft (1) +c& (2) , and examine the terms of the second order of small- ness. The equation with k = n gives E ®>C <°> = S' V r CD whence E, « (2) =2 ' \V I 2 mn\ E <®—E <°> (38.9) (we have substituted cj" = V m J(E n W-E m ^) t and used the fact that, since the operator V is Hermitian, V mn = V nm *). We notice that the correction in the second approximation to the energy of the normal state is always negative; for, since £ n <°> then corresponds to the lowest value of the energy, all the terms in the sum (38.9) are negative. The further approximations can be calculated in an exactly similar manner. The results obtained can be generalised at once to the case where the operator H has also a continuous spectrum (but the perturbation is applied, as before, to a state of the discrete spectrum). To do so, we need only add to the sums over the discrete spectrum the corresponding integrals over the continuous spectrum. We shall distinguish the various states of the continu- ous spectrum by the suffix v, which takes a continuous range of values; by v we conventionally understand an assembly of values of quantities sufficient for a complete description of the state (if the states of the continuous spec- trum are degenerate, which is almost always the case, the value of the energy alone does not suffice to determine the state).f Then, for instance, we must write instead of (38.7) and similarly for the other formulae. It is useful to note also the formula for the perturbed value of the matrix element of a physical quantity/, calculated as far as terms of the first order by using the functions n = ^<°>+0 n tt> f with ^a) g i ven by (38.7). The t Here the wave functions «£„(») must be normalised by delta functions of the quantities v. 132 Perturbation Theory §38 following expression is easily obtained: ^' *W*m (0) ^' V km f nk W f =f (o)+ > u > (38 U) k *-"n. ^k k ^m ^k In the first sum k ^ n, while in the second k ^ m. PROBLEMS Problem 1. Determine the correction ^„ (2) in the second approximation to the eigen- functions. Solution. The coefficients Ck (2) (k ^ n) are calculated from equations (38.4) with k ^ n, written out up to terms of the second order, and the coefficient c„< 2 > is chosen so that the function t(i n = ^n (0) +^n (1) +^» (2) is normalised up to terms of the second order. As a result we find *•* =22 j£S^-*-2 ^---**.-2'fe^ m k ul nk w nm m '* w nm m n w nm where we have introduced the frequencies Problem 2. Determine the correction in the third approximation to the eigenvalues of the energy. Solution. Writing out the terms of the third order of smallness in equation (38.4) with k — n, we obtain „ (3) _ V"V' ^nmYmkVkn NT^' \Vn, Problem 3. Determine the energy levels of an anharmonic linear oscillator whose Hamil- tonian is Solution. The matrix elements of x 3 and ** can be obtained directly according to the rule of matrix multiplication, using the expression (23.4) for the matrix elements of x. We find for the matrix elements of x 3 that are not zero (^)n- 3 ,« = («■)..-. = (hlmcof/W[Mn-l)(n-2)], (*■)-!.. = (^)«,«-i = (W 3/ V(M8). The diagonal elements in this matrix vanish, so that the correction in the first approximation due to the term ax 3 in the Hamiltonian (regarded as a perturbation of the harmonic oscillator) is zero. The correction in the second approximation due to this term is of the same order as that in the first approximation due to the term fa*. The diagonal matrix elements of a: 4 are (**)«,« = (h/mojy . |(2» 2 +2 W +1). Using the general formulae (38.6) and (38.9), we find the following approximate expression for the energy levels of the anharmonic oscillator: 15 a 2 / h \ V 11\ 3 / h \ 2 E n =ha>{n+l)--—{ — ) (»■+„+_ )+-p(_) < w2+w +*)- 4 hat \mco/ \ 30/ 2 \mco/ §39 The secular equation 133 §39. The secular equation Let us now turn to the case where the unperturbed operator # has de- generate eigenvalues. We denote by </r n <°>, ^ n ,<o) f ... the eigenfunctions be- longing to the same eigenvalue £ n <°> of the energy. The choice of these func- tions is, as we know, not unique; instead of them we can choose any s (where s is the degree of degeneracy of the level £ n <°>) independent linear combina- tions of these functions. The choice ceases to be arbitrary, however, if we subject the wave functions to the requirement that the change in them' under the action of the small applied perturbation should be small. At present we shall understand by if, n «>\ f n S<>\ ... some arbitrarily selected unperturbed eigenfunctions. The correct functions in the zeroth approxima- tion are linear combinations of the form %<°ty n <°> + c n > «»ip n ' «»+.... The co- efficients in these combinations are determined, together with the corrections in the first approximation to the eigenvalues, as follows. We write out equations (38.4) with £=»,»',..., and substitute in them, in the first approximation, E = £ n «» + £(i) ; for the quantities c h it suffices to take the zero-order values c n = c n <°\ c n . = cj°\ ... ; c m = for m # n, n', ... . We then obtain or S(F nn ,-#»S wn 0^<°>=0, (39.1) where n, n' take all values denumerating states belonging to the given un- perturbed eigenvalue £ TC <°>. This system of homogeneous linear equations for the quantities c„< 0) has solutions which are not all zero if the determinant of the coefficients of the unknowns vanishes. Thus we obtain the equation |F nn ,-2?a>S nn ,|=(). (39.2) This equation is of the rth degree in E™ and has, in general, s different real roots. These roots are the required corrections to the eigenvalues in the first approximation. Equation (39.2) is called the secular equation.} We notice that the sum of its roots is equal to the sum. of the diagonal matrix elements V nn > Vnn; »• ( tni s being the coefficient of [2?fl)]*-i in the equation). Substituting in turn the roots of equation (39.2) in the system (39.1) and solving, we find the coefficients <: n < > and so determine the eigenfunctions in the zeroth approximation. We notice that, as a result of the perturbation, an originally degenerate energy level ceases in general to be degenerate (the roots of equation (39.2) are in general distinct); the perturbation removes the degeneracy, as we say. The removal of the degeneracy may be either total or partial (in the latter case, after the perturbation has been applied, there remains a degeneracy of degree less than the original one). t The name is taken from celestial mechanics. 134 Perturbation Theory §39 It may happen that all the matrix elements for transitions between the states «, n\ ... with a single energy are zero. The correction to the energy then vanishes in the first approximation. Let us calculate the correction in the second approximation for this case. In equation (38.4) with k = n we put on the left-hand side E = E n w +E®\ and write c n w in place of c n . Only the terms with m ^ w, »', ... on the right-hand side are different from zero, and since c m w = we have EPcf* = S V nm c^\ (39.3) m The equation (38.4) with k = m ^ n, ri , ... , on the other hand, gives, correct to terms of the first order, n whence c 0> = V Vmn ' c .<o> n Substituting in (39.3), we obtain £(2) c (0) = V c ,(0)VM!!!l n Z-, n Z-, e <°>— £• <°> ' This system of equations for the c n (0) now replaces the system (39.1); the condition that these equations are compatible is xr-^ V V > ^ nm mn -&% nn . £i F. (o)— E (°)— E <°> = 0. (39.4) Thus here also the corrections to the energy are calculated as the roots of a secular equation, in which, instead of the matrix elements V nn >, we now have the sums ' 2* nm* mn' E (°)— E (o) PROBLEMS Problem 1. Determine the corrections to the eigenvalue in the first approximation and the correct functions in the zeroth approximation, for a doubly degenerate level. Solution. Equation (39.2) here has the form V u -E»> V 12 V» V,,-E^ = (the suffixes 1 and 2 correspond to two arbitrarily chosen unperturbed eigenfunctions i/^ * §39 The secular equation 135 and 0a (o) of the doubly degenerate level in question). Solving, we find £ (1) = *[(F U + v 22 )±v{(v 11 - v 22 y+4\ F 12 p}]. (i) Solving also equations (39.1) with these values of £<D, we obtain for the coefficients in the correct normalised function in the zeroth approximation, i/r«» = ci< >«Ai<°) + c2< ><A2<°> the values ' w= f v 12 r . F,i-F« 1U b r c a <°> fl± ^^ I)* |L V{(^i-^ 2 ) 2 +4|F 12 |2}J/ , (2) - =fc l^LTlT ^^ I)* (2| V 12 \ I V{(V X1 - V 22 )*+4\ F 12 p}Jj ' Problem 2. Derive the formulae for the correction to the eigenfunctions in the first approximation and to the eigenvalues in the second approximation. Solution. We shall suppose that the correct functions in the zeroth approximation are chosen as the functions *„(»>. The matrix V nn , defined with respect to these is clearly diagonal with respect to the suffixes n, n' (belonging to the same group of functions of a degenerate Zfx i*^ ■ di f go ? al elements V-> Vww are equal to the corresponding corrections *V \ i4n»* ', ... in the first approximation. Let us consider a perturbation of the eigenfunction if> n M, so that in the zeroth approxima- tion E = E n (0) , c„(°) = 1, c »(°) = for m # n. In the first approximation E = E n ^ + V nn c » - 1 + c » (1, » Cr» = c m M. We write out from the system (38.4) the equation with k ^ n n' retaining in it terms of the first order: ' ''"' W-E^p = V kn c n «» = V kn , whence <* (1) = VjJVZjn-EP) for h ± n, n\ ... . (1) Next we write out the equation with k = n\ retaining in it terms of the second order: E n <»c n 'to = V n > n > c„'<«-f E' V > c « (the terms with m = n, n', ... are omitted in the sum over m). Substituting E <*> = V and the expression (1) for c m W, we obtain for n' # n " "" 1 xrV V > V (X) ~ X ' n m r mn V — V • > ^-i F (0)_ F (0) ' (2) (In this approximation the coefficient c»U> is zero.) Formulae (1) and (2) determine the correction 0.W = Sc^tyJ") to the eigenfunctions in the first approximation. Finally, writing out the second-order terms in equation (38.4) with k = n, we obtain for the second-order corrections to the energy the formula V V ' nm v mn 2) = y _Vnm_ r E « (0) - -£<<»' (3) which is formally identical with (38.9). Problem 3. At the initial instant * = 0, a system is in a state &<») which belongs to a doubly degenerate level. Determine the probability that, at a subsequent instant t, the system will be in the state &«» with the same energy; the transition occurs under the action of a constant perturbation. Solution. We form the correct functions in the zeroth approximation, 136 Perturbation Theory §40 where c lt c t ; c^, c t ' are two pairs of coefficients determined by formulae (2) of Problem 1 (for brevity, we omit the index (°) on all quantities). Conversely, «Ai = ; — ■ The functions ifi and ifi' belong to states with perturbed energies E+EW and E+EW, where EW and EW are the two values of the correction (1) in Problem 1. On introducing the time factors we pass to the time-dependent wave functions : g-dfldEt T x = ; —[c^e-QlMEt-Citft'e-QlME <] tjtg — ^l ^2 (at time t = 0, Tj = fa). Finally, again expressing tft, tfi' in terms of fa, fa, we obtain Yj as a linear combination of fa and fa, with coefficients depending on time. The squared modu- lus of the coefficient of fa determines the required transition probability w 12 . Calculation gives / Q \ e -(i/h)E Kl 't_ e -(i/h)E Kl ' <|2 j or, substituting formulae (1) and (2) from Problem 1, 2|r M |» MV 12 \'+(V 11 -V 22 f l-cos^-VliV^-V^f+^V^ty^ . (1) We see that the probability varies periodically with time, with frequency (EW—EW)jH. For times t which are small compared with the period in question, the expression in the braces, and therefore W12, is proportional to t 2 : W12 = \Vi2\ 2 t 2 /H z . This formula can be very simply obtained by the method given in the next section (using equation (40.4)). §40. Perturbations depending on time Let us now go on to study perturbations depending explicitly on time. We cannot speak in this case of corrections to the eigenvalues, since, when the Hamiltonian is time-dependent (as will be the perturbed operator & = $ Q + + !?"(£)), the energy is not conserved, so that there are no stationary states. The problem here consists in approximately calculating the wave functions from those of the stationary states of the unperturbed system. To do this, we shall apply a method analogous to the well-known method of varying the constants to solve linear differential equations (P. A. M. Dirac 1926). Let Y& (0) be the wave functions (including the time factor) of the stationary states of the unperturbed system. Then an arbitrary solution of the unperturbed wave equation can be written in the form of a sum: T = 2flfc¥V°). We shall now seek the solution of the perturbed equation ih dW/dt = (# + P)¥ (40.1) in the form of a sum T = S<z fc (0T fc <o>, (40.2) where the expansion coefficients are functions of time. Substituting (40.2) §40 Perturbations depending on time 137 in (40.1), and recalling that the functions Y^ satisfy the equation we obtain k "t k Multiplying both sides of this equation on the left by T OT (0) * and integrating, we have where are the matrix elements of the perturbation, including the time factor (and it must be borne in mind that, when V depends explicitly on time, the quanti- ties V mk also are functions of time). As the unperturbed wave function we take the wave function of the nth stationary state, for which the corresponding values of the coefficients in (40.2) are a n (0) = 1, a k w = for k # n. To find the first approximation, we seek a h in the form a k = a k w +a k a) , substituting a k = a k {0) on the right-hand side of equation (40.3), which already contains the small quantities V mk . This gives ihda k V>ldt = V kn (t)- (40.4) In order to show the unperturbed function to which the correction is being calculated, we introduce a second suffix in the coefficients a k , writing Y n = Sa fcn (*)V. Accordingly, we write the result of integrating equation (40.4) in the form a*» (1) = -(«/*) J* V kn (t) & = -(ilh) j V kn e*»*nt dt, (40.5) where we have introduced the frequencies (o kn = (E k w —E n {0) )lh. This determines the wave functions in the first approximation. We can similarly determine the subsequent approximations (in practice, however, the first approximation is adequate in the majority of cases). Let us now consider in more detail the case of a perturbation which is periodic with respect to time, of the form ■p z= p e -i<ot + @ e ic*t i (40.6) 138 Perturbation Theory §40 where P and are operators independent of time. Since V is Hermitian, we must have V nm = V mn *, or J? p-ioit 1/7 oittit _ K 1 * p i(oiir; # p -io>t from which which determines the relation between the operators and P. Using this relation, we have = Ftoe* a *- ai *+F ia *e 1 l a> *»+ ai *. (40.8) Substituting in (40.5) and integrating, we obtain the following expression for the expansion coefficients of the wave functions : Fto/to**-"* F nl *e*- M kn+<»V K(<x) kn —aj) H(aj kn +co) These expressions are applicable if none of the denominators vanishes,! i.e. if for all k (and the given n) EjP-E® * ±hco. (40.10) In a number of applications it is useful to have expressions for the matrix elements of an arbitrary quantity /, defined with respect to the perturbed wave functions. In the first approximation we have JnmSJ) = Jnm \f)ijnm \r)i where fnm ( °Kt) = jT n W*/T w (0)d ? =/ B J¥M /™ (1) (0 = j W 0) */ V+^« (1) */^ 0) ] dq. Substituting here T n < 1) = S a &n cl,, *V 0) » with « fcTC (1) determined by formula (40.9), it is easy to obtain the required expression fnnP(t) = -*"*"* V [\ f<*® F *» + f*»® F ** 1 e -^< + ^ lL^(a) fcm — cu) ^(eo^+a^J + r /A* A^V-i^a (40 . n) L^(co ftTO +co) n{oi kn — u))j ) f More precisely, if none is so small that the quantities aje n W are no longer small compared with unity. §40 Perturbations depending on time 139 This formula is applicable if none of its terms becomes large, i.e. if none of the frequencies o)^ n , co^ m is too close to co. For <o = we return to formula (38.11). In all the formulae given here, it is understood that there is only a discrete spectrum of unperturbed energy levels. However, these formulae can be immediately generalised to the case where there is also a continuous spectrum (as before, we are concerned with the perturbation of states of the discrete spectrum) ; this is done by simply adding to the sums over the levels of the discrete spectrum the corresponding integrals over the continuous spectrum. Here it is necessary for the denominators a) kn ±eo in formulae (40.9), (40.11) to be non-zero when the energy E k w takes all values, not only of the discrete but also of the continuous spectrum. If, as usually happens, the continuous spectrum lies above all the levels of the discrete spectrum, then, for instance, the condition (40.10) must be supplemented by the condition EnJQ-E® > ha>, (40.12) where J? min (0) is the energy of the lowest level of the continuous spectrum. PROBLEM Determine the change in the nth and mth solutions of Schrodinger's equation in the presence of a periodic perturbation (of the form (40.6)), of frequency a» such that £' m < >— EJ'1 = h(a>+e), where e is a small quantity. Solution. The method developed in the text is here inapplicable, since the coefficient aj 1 ' in (40.9) becomes large. We start afresh from the exact equations (40.3), with V mlc (t) given by (40.8). It is evident that the most important effect is due to those terms, in the sums on the right-hand side of equations (40.3), in which the time dependence is determined by the small frequency to mn — to. Omitting all other terms, we obtain a system of two equa- tions : ihdajdt = F mn e*<° mn -<»)ta n = F mn e™a ni ihdajdt = F mn *e-^a m . We make the substitution a e iet = h and obtain the equations ihd m = F mn b nt th(b n —t€b n ) = F mn *a m . Eliminating a m , we have h n -kb n +\F mn \*b n l¥ = 0. We can take as two independent solutions of these equations a n = AeW, a m = ~Ah^ x e^\F m * (1) and a n = Be-W, a m = Bkx^-W/F^*, (2) 140 Perturbation Theory §41 where A and B are constants (which have to be determined from the normalisation condition), and we have used the notation «1 = -h+Vih 2 +\Fmn\ 2 m a 2 =h+V[k 2 +\Fmn\Wl Thus, under the action of the perturbation, the functions Y„(°>, T m <°) become a n ¥ n <°> + +a m Y TO <°), with a n and a m given by (1) and (2). Let the system be in the state T m <°> at the initial instant (t — 0). The state of the system at subsequent instants is given by a linear combination of the two functions which we have obtained, which becomes Y ro <°) for t — 0: I*" win 1 ■ — e-^/2sinV[ie 2 +|F mn |2/^ 2 ]« T n «». (3) The squared modulus of the coefficient of Y n (°) is ^•\r mr> /* 2 e 2 +4|^„ ^l-cosV[ € 2 +4|^ wn | 2 /^]0. (4) This gives the probability of finding the system in the state VF n (°> at time t. We see that it is a periodic function with period 2irh(e 2 h 2 +4\F mn \ z )~ i , and varies from to 4l.F mB |V(J* 2 e 2 4- +4| J F ran | 2 ). For e = (exact resonance) the probability (4) becomes *(l-cos2|F mn |^). It varies periodically, with period nH/\F mn \, between and 1; in other words, the system makes periodic transitions from the state T m <°) to the state Y,,* ). §41. Transitions under a perturbation acting for a finite time Let us suppose that the perturbation V(t) acts only during some finite interval of time (or that V(t) diminishes sufficiently rapidly as t -> ± oo). Let the system be in the nth stationary state (of a discrete spectrum) before the perturbation begins to act (or in the limit as t -> — oo). At any subsequent instant the state of the system will be determined by the function T = 2 a k1 ^¥ k w , where, in the first approximation, «*n = a k v> = -- f V^e^Jdt for k*n, a nn = l+a Bn <« = 1-- I V n (41.1) d*; the limits of integration in (40.5) are taken so that, as t -> — oo, all the tf &n (1> tend to zero. After the perturbation has ceased to act (or in the limit t -> + oo), the coefficients a hn take constant values a kn {co), and the system is in §41 Transitions under a perturbation acting for a finite time 141 the state with wave function Y = S« ftn (a>W 0) , which again satisfies the unperturbed wave equation, but is different from the original function ¥ n < 0) . According to the general rule, the squared modulus of the coefficient a kn (co) determines the probability for the system to have an energy E k <°\ i.e. to be in the Mi stationary state. Thus, under the action of the perturbation, the system may pass from its initial stationary state to any other. The probability of a transition from the original (wth) to the Ath stationary state is h*\J V*^°>*f dt (41.2) Let us now consider a perturbation which, once having begun, continues to act for an indefinite time (always, of course, remaining small). In other words, V(t ) tends to zero as t -» — oo and to a finite non-zero limit as t -> + oo. Formula (41.2) cannot be applied directly here, since the integral in it diverges. This divergence, however, is physically unimportant and can easily be removed. To do this, we integrate by parts : -— IJ^*--[^T+J— — *• The value of the first term vanishes at the lower limit, while at the upper limit it is formally identical with the expansion coefficients in formula (38.7); the presence of an additional periodic factor e %0iknt is merely due to the fact that the a kn are the expansion coefficients of the complete wave function T, while the c kn in §38 are the expansion coefficients of the time-independent function *Jj. Hence it is clear that its limit as t -*■ oo gives simply the change in the original wave function *F n (0 > under the action of the "constant part" F( + oo) of the perturbation, and consequently has no relation to transitions into other states. The probability of a transition is given by the squared modulus of the second term and is w n k = — **"*■' dt W I J dt / (41.3) The derivation is also valid when the transition is from a state of the discrete spectrum to a state of the continuous spectrum. The only difference is that here we have the probability of the transition from a given (nth) state to states in a range of values of v (see §38) from v to v + dv, so that, for example, formula (41.2) must be written °° 1 I /* 2 dw nv = — Vv n e i( °vnt dt dv. (41.4) h 2 \ J 142 Perturbation Theory §41 If the perturbation V(t) varies little during time intervals of the order of the period l/co fcn the value of the integral in (41.2) or (41.3) will be very small. In the limit when the applied perturbation varies arbitrarily slowly, the probability of any transition with change of energy (i.e. with a non-zero frequency ca kn ) tends to zero. Thus, when the applied perturbation changes sufficiently slowly {adiabatically), a system in any non-degenerate stationary state will remain in that state (see also §53). In the opposite limiting case of a very rapid, "instantaneous" application of the perturbation, the derivatives dVjcn/dt become infinite at the "instant of application". In the integral of (8V kn ldt)e i ' ^ t , we can take outside the integral the comparatively slowly varying factor e* w *»* and use its value at this instant. The integral is then found at once, and we obtain v>nk=\V kn \ 2 /h*a> kn *. (41.5) The transition probabilities in instantaneous perturbations can also be found in cases where the perturbation is not small. Let the system be in a state described by one of the eigenfunctions ifj n (Q) of the original Hamiltonian Ho. If the change in the Hamiltonian occurs instantaneously (i.e. in a time short compared with the periods l/to to of transitions from the given state n to other states), then the wave function of the system is "unable" to vary and remains the same as before the perturbation. It will no longer, however, be an eigenfunction of the new Hamiltonian H of the system, i.e. the state j/f w <°> will not be a stationary state. The probabilities w nk for transitions of the system into the new stationary states are determined, according to the general rules of quantum mechanics, by the coefficients in the expansion of the function ift n (°> in terms of the eigenfunctions *p n of the Hamiltonian i?: ™nk = | \ W m H* <*/. (41.6) We shall show how this general formula becomes (41.5) if the change V = A— i? in the Hamiltonian is small. We multiply the equations by ifjjc* and i/%<°)* respectively, integrate with respect to q and subtract. Using also the self-conjugacy of the operator ]2, we obtain (E k -E n ™) j <A**<An (0) d? = j 0**Jty» (O) 6q. If the perturbation V is small, in the first approximation we can replace Ejc by the adjoining unperturbed level E^, and the wave function ifj^ (on the right-hand side of the equation) by the corresponding function i/jjc^. This gives f «A*W 0) dq = [ W<»*fyn {0) dq, J natjcn J and formula (41.6) becomes (41.5). §41 Transitions under a perturbation acting for a finite time 143 PROBLEMS Problem 1. A uniform electric field is suddenly applied to a charged oscillator in the ground state. Determine the probabilities of transitions of the oscillator to excited states under the action of this perturbation. Solution. The potential energy of the oscillator in the uniform field (which exerts a force F on it) is U(x) = \moi 2 x 2 —Fx = %maj 2 (x—xo) 2 + constant (where xo = F/mw 2 ), i.e. has still the pure oscillator form but with the equilibrium position shifted. Hence the wave functions of the stationary states of the perturbed oscillator are <lik(x— xo), where >/ijc(x) are the oscillator functions (23.12); the initial wave function is tfio(x) (23.13). Using these functions and the expression (23.11) for the Hermite polynomials, we find 00 00 f (-1)* r d * «£o (0 ty* dx = — — e-£o<>/2 e -Uo e -^+2Ko d£, J y/(2*irk\) J d£* -00 v \ / _ K) with the notation |o = xo-\Z(tnu)jh). On integrating k times by parts, the integral on the right becomes 00 —00 Thus the transition probability (41.6) is £ 2* As a function of the number k it represents a Poisson distribution for which the mean value of k is k = #„ z = F*l2mha>*. Perturbation theory is applicable when F is small, so that k <^ 1. Then the excitation probabilities are small, and decrease rapidly with increasing k. The largest is woi = k. In the opposite case of large F (k ^> 1), excitation of the oscillator occurs with very high probability: the probability that the oscillator will remain in the normal state is woo = e~*. Problem 2. The nucleus of an atom in the normal state receives an impulse which gives it a velocity v; the duration t of the impulse is assumed short in comparison both with the electron periods and with ajv, where a is the dimension of the atom. Determine the probability of excitation of the atom under the influence of such a "jolt" (A. B. Migdal 1939). Solution. We use a frame of reference K' moving with the nucleus after the impact. By virtue of the condition r <^ afv, the nucleus may be regarded as practically stationary during the impact, so that the co-ordinates of the electrons in K' and in the original frame K immediately after the perturbation are the same. The initial wave function in K' is *i>o = ^0 exp( — tq . 2 r fl ), q = mvjh, where ^o is the wave function of the normal state with the nucleus at rest, and the summation 144 Perturbation Theory §41 in the exponent is over all Z electrons in the atom; see §15, Problem 2. The required prob- ability of transition to the Ath excited state is now given, according to (41.6), by WOK J fa* exp(-iq . 2 r a )^ dFi ... dV z \2 . In particular, if qa <^ 1 , then by expanding the exponential factor in the integrand and noting that the integral of <pk*<jio is zero because the functions ifio and ipk are orthogonal, we obtain »0* | f^»(q.Sro)0odFi...dKz| 2 . J a Problem 3. Determine the total probability of excitation and ionisation of an atom of hydrogen which receives a sudden "jolt" (see Problem 2). Solution. The required probability can be calculated as the difference 1 -WOO = 1 - 1 J Wer*** d j/|2 f where woo is the probability that the atom will remain in the ground state (^o = TT~ U2 e~ r l a being the wave function of the ground state of the hydrogen atom, with a the Bohr radius). Calculation of the integral gives l-«oo= l-l/(l+& 2 « 2 ) 4 . In the limiting case qa <^ 1 this probability tends to zero as q 2 a 2 , while for qa ^> 1 it tends to unity as 1 — (2/qa) s . Problem 4. Determine the probability that an electron will leave the K-shell of an atom with large atomic number Z when the nucleus undergoes /3-decay. The velocity of the ^-particle is assumed large in comparison with that of the iC-electron (A. B. Migdal and E. L. Feinberg 1941). Solution.! In the conditions stated the time taken by the ^-particle to pass through the iC-shell is small compared with the period of revolution of the electron, so that the change in the nuclear charge can be regarded as instantaneous. The perturbation is here represented by the change V = 1/r in the field of the nucleus when the change in its charge is small (1 compared with Z). According to (41.5) the transition probability for one of the two iC-shell electrons with energy Eo = —\Z 2 (here and below we use the fact that the state of the K- electrons is hydrogen-like; see §74) to a state of the continuous spectrum with energy E = $k 2 in the range dE = k dk is dw = 2 d&. (£2 + Z2)2 In the range which determines the matrix element Voic, the important part is that of short distances (~l/Z) from the nucleus, in which the hydrogen-like expression can again be used for the wave function of a state of the continuous spectrum. The final state of the electron must have angular momentum I = (the same as that of the initial state). By means of the functions Rio and i?&o (normalised on the k scale) derived in §36 and formula (f.3) in the Mathematical Appendices we find J 1 \ 4 Vk ( 1 + tkjZ)W*( 1 - ikjZ)-iz/k \r/ok Ok y/(l-er*»V*) 1+&2/Z2 t In Problems 4 and 5, atomic units are used. | In the calculation it is convenient to use Coulomb units and then return to atomic units in the final result. §41 Transitions under a perturbation acting for a finite time 145 and, since |(l+w)</«| 2 = exp[-(2/a)tan- 1 a], we obtain finally dw = f(k/Z)k dk, with 1 /(°0 = : TT ex P[~(4/a) tan" 1 a]. The limiting values of the function /(a) are e~4 for a <^ 1 and v.\2-n for a > 1. The total probability of ionisation of the X-shell is obtained by integration of dw over all energies of the emergent electron. A numerical evaluation gives w = 0-652T 2 . Problem 5. Determine the probability of emergence of an electron from the i^-shell of an atom with large Z in oc-decay of the nucleus. The velocity of the oc-particle is small compared with that of the ^-electron, but the time which it takes to leave the nucleus is small in comparison with the time of revolution of the electron.! Solution. After the emergence of the oc-particle, the perturbation acting on the electron is adiabatic. The required effect is therefore determined essentially by the interval of time close to the "instant of application' ' of the perturbation which destroys the adiabaticity, when the a-particle, leaving the nucleus and moving freely, is still at a distance small compared with the radius of the iC-orbit. The perturbation V which causes the ionisation of the atom is here represented by the deviation of the combined field of the nucleus and the a-particle from the purely Coulomb field Zjr. The dipole moment of two particles with atomic weights 4 and A—\, and charges 2 and Z— 2, at a distance vt apart (where v is the relative velocity of the nucleus and the a-particle), is{ 2(4 - 4) -(Z- 2)4 2(A-2Z) vt = vt. A A Hence the dipole term in the field of the nucleus and the a-particle is|| 2(A-2Z) z V = vt—, A f3' where the 2-axis is in the direction of the velocity v. The matrix element of this perturbation reduces to that of z: taking the matrix element of the equation of motion of the electron z = —Zz/r 3 , we obtain (zlr*)o k = (E-E )*z iclZ. The required transition probability for one of the two electrons in the K-shell is, by (41.2), dw = 2| J Voke^Eo-Eu dt\ dk o $(A-2Z) 2 v2 A 2 Z* Fo*| 2 dk; t This problem was first discussed by A. B. Migdal (1941). } The necessity of simultaneously considering the motion of the a-particle and that of the nucleus in this problem has been pointed out by J. S. Levinger (1953). || If the difference A— 2Z is small, it may be necessary to take account of the next (quadrupole) term also. 146 Perturbation Theory §42 to calculate the integral, we include in the integrand an additional damping factor e~ M with A > 0, and then make A -> in the result. To calculate the matrix element of z — r cos 6, we note that, since the orbital angular momentum in the initial state is I = 0, cos has a non- zero matrix element only for the transition to a state with 1=1, and |(cos0)oi| 2 = (COS0) OO = | and M 2 = i\r 0k \ 2 . Calculating rok by means of the radial functions Roo and Rki, we find 211(A-2ZW dzv = f(k/Z)k dk, 3A 2 Z%l+k*IZ 2 )S the function / being as in Problem 4. §42. Transitions under the action of a periodic perturbation The results are different for the probability of transitions to the states of the continuous spectrum under the action of a periodic perturbation. Let us suppose that, at some initial instant t = 0, the system is in the wth station- ary state of the discrete spectrum. We shall assume that the frequency co of the periodic perturbation is such that hco>E min -E^\ (42.1) where E min is the value of the energy where the continuous spectrum begins. It is evident from the results of §40 that the chief part will be played by states with energies E v very close to the resonance energy E n w +hco, i.e. those for which the difference co vn —co is small.f For this reason it is sufficient to consider, in the matrix elements (40.8) of the perturbation, only the first term (with the frequency co vn — co close to zero). Substituting this term in (40.5) and integrating, we obtain " vn = -t K„(t) dt = -F- _. (42.2) hj % n -w) The lower limit of integration is chosen so that a vn = for t =0, in accord- ance with the initial condition imposed. Hence we find for the squared modulus of a vn KJ 2 = l*UI 2 • 4 sin*[Kc^-a,)*]//*V,,„-") 2 . (42.3) It is easy to see that, for large t, this function can be regarded as propor- tional to t. To show this, we notice that sin 2 a£ lim = 8(a). (42.4) f->oo Trta? For when a # this limit is zero, while for a = we have (sin 2 otf)/ta 2 = t, so that the limit is infinite; finally, integrating over a from — oo to +a>, f We recall that the suffix v refers to the continuous spectrum (see the end of §38). §™ Transitions in the continuous spectrum 147 we have (with the substitution at = £) 1 r sin 2 af 1 j* sin 2 £ 1 r sm 2 xt 1 f sin 2 £ - — —da =- d$ = 1. 77-J to 2 77 J £ 2 Thus the function on the left-hand side of equation (42.4) in fact satisfies all the conditions which define the delta function. Accordingly, we can write for large t kJ 2 = ( i/£VJ 2 ^s(K«-K), or, substituting hco vn = E v -E n ™ an d using the fact that 8(ax) = (l/a)8(«), \a vn \* = (2*lh)\FJWB v -E n m-ha>)t. The expression |a m | 2 dv is the probability of a transition from the original state to one in the interval from v to v+dv (cf. §5). We see that, for large t, it is proportional to the time interval elapsed since t = 0. The probability dzv nv of the transition per unit time isf d«v = (2iT/h)\FJ*8(E,,-EM-hco) dv. (42.5) As we should expect, it is zero except for transitions to states with energy E v = E n W+hco. If the energy levels of the continuous spectrum are not degenerate, so that v can be taken as the value of the energy alone, then the whole "interval" of states dv reduces to a single state with energy E = E n w + + hm, and the probability of a transition to this state is VnE = {27rlh)\F En \\ (42.6) §43. Transitions in the continuous spectrum One of the most important applications of perturbation theory is to calculate the probability of a transition in the continuous spectrum under the action of a constant (time-independent) perturbation. We have already mentioned that the states of the continuous spectrum are almost always degenerate. Having chosen in some manner the set of unperturbed wave functions cor- responding to some given energy level, we can put the problem as follows. It is known that, at the initial instant, the system is in one of these states; it is required to determine the probability of the transition to another state with the same energy. If we denote the initial state by the suffix v , then for transitions to states between v and v+dv we have at once from (42.5) (putting co = and changing the notation) <K.„ = ( 27r ^)l V w} 2 KE-E v ) dv. (43.1) This expression is, as we should expect, zero except for E V =E V : under .*!•«? iS i CaSy t0 ™ rify th&t l °? tBking acCOunt of the second term in < 40 - 8 >' which we have omitted, additional expressions are obtained which, on being divided by t, tend to zero as *-» +00. 148 Perturbation Theory §43 the action of a constant perturbation, transitions occur only between states with the same energy. It must be noticed that, for transitions from states of the continuous spectrum, the quantity dw VtV cannot be regarded directly as the transition probability; it is not even of the right dimensions (1/time). Formula (43.1) represents the number of transitions per unit time, and its dimensions depend on the chosen method of normalisation of the wave functions of the continuous spectrum. Let us calculate the perturbed wave function which at the initial instant coincides with the initial unperturbed function $ Vt w . According to formula (42.2) (putting co = and changing the notation) we have AilME v -E )t_\ a a) = V **HM w It'll "• E vr E v The perturbed wave function has the form T„ o =T v W+J^T/)dv, or r r \- e (iimE Vo -E v )t -, 1 J Ev ~ Ev J (43.2) where the integration is extended over the whole continuous spectrum.f Let us ascertain the limiting form of this function for large t. To do so, we separate from dv the differential dE v of the energy (writing dv = dE v dr, where dr is the product of the differentials of the remaining quantities which determine a state in the continuous spectrum), and formally regard E v as a complex variable. The integral f \— e {i/ME Vo -E v )t r -*' m E..-E.. "* in (43.2) is taken along the real axis. We slightly displace the path of inte- gration into the lower half-plane; this can be done without changing the value of the integral, since the integrand has no singularities on the real axis. The integral can then be divided into two parts r dE r Vw <A V (0) J vv ^ v E v -E v J E v -E v (these integrals have no meaning when the integration is along the real axis, since they diverge at the point E v = E v ). Since im E v < on the contour of integration, the second of the above integrals tends to zero as t tends J:o t If there is also a discrete spectrum, then we must add to the integral in this formula (and subse- quent ones) the appropriate sum over the states of the discrete spectrum. §43 Transitions in the continuous spectrum 149 infinity (because of the factor exp[fr-iim(£ w )*] in the integrand). In the first integral, we can again make the path of integration the real axis, but pass round the point E v = E v ^ below. This method of integration can be conven- iently represented in another form by adding to the constant E„ in the denominator of the integrand a small positive imaginary quantity tS. The pole of the integrand is thereby shifted into the upper half-plane, and the integration can be taken simply along the real axis (which now passes below the pole), after which 8 is allowed to tend to zero. Thus we obtain for the wave function the expression T ». = [fv t m + j E "° E +i8 W 0) dJje-UmE^ S ^o+ . (43.3) The time factor shows that this function belongs, as it should, to the same energy E Vt as the initial unperturbed function. In other words, the function K =K {0) + » — ^ (0)d » satisfies Schrodinger's equation (# + P)^ = E v $ n . For this reason it is natural that the expression above should correspond exactly to formula (38.7).f The calculations given above correspond to the first approximation of perturbation theory. It is not difficult to calculate the second approximation as well. To do this, we must derive the formula for the next approximation to Y Vo ; this is easily effected by using the method of §38 (now that we know the method of dealing with the "divergent" integrals). A simple calculation gives the formula with the same method of indentation in the integrals. Comparing this expression with formula (43.3), we can write down the corresponding formula for the probability (or, more precisely, the number) of transitions, by direct analogy with (43.1): dzo = — n r V VV .V V . V , 2 Kv+ F -p dv ' 8 ^.~^) d". (43.5) It may happen that the matrix element V vv% for the transition considered >u f Sta iS ng fr u m l he latter formuIa > the wa y ^ which the integral must be taken can be found from !»L C ^f? 0!? " *? "J"** *? ^Pression for ^. at large distances must contain only an outgoing (and not an ingoing) wave (see §134). " i s"«'s 150 Perturbation Theory §44 vanishes. The effect is then zero in the first approximation, and we have for the number of transitions ^ J i\i^ d ^- E ^- (43 ' 6) In applications of this formula, the point E v , = E Va is not usually a singularity of the integrand; the manner of integrating with respect to E v , is therefore unimportant in general, and the integral can be taken along the real axis. The states v for which V vv , and V v . Vv are not zero are usually called intermediate states for the transition v -> v. It may happen that the transi- tion v -> v can take place not through one but only through several successive intermediate states. Formula (43.6) can be at once generalised to such cases. Thus, if two intermediate states are needed, we have 2tt\ r r VwnVv.'v.Vv^, 7/ 2 dw = — I I dv dv "•" h\)](E v -E v .){E-E v .) 8(E-E Vo )dv. (43.7) §44. The uncertainty relation for energy Let us consider a system composed of two weakly interacting parts. We suppose that it is known that at some instant these parts have definite values of the energy, which we denote by E and e respectively. Let the energy be measured again after some time interval At; the values E', e obtained are in general different from E, e. It is easy to determine the order of magnitude of the most probable value of the difference E' + e —E— e which is found as a result of the measurement. According to formula (42.3) with <o = 0, the probability of a transition of the system (after time t), under the action of a time-independent perturbation, from a state with energy E to one with energy E' is proportional to sm*[(E'-E)tl2h]l(E'-E)\ Hence we see that the most probable value of the difference E'—E is of the order of hjt. Applying this result to the case we are considering (the perturbation being the interaction between the parts of the system), we obtain the relation \E+ € -E'-€'\At~h. (44.1) Thus the smaller the time interval A*, the greater the energy change that is observed. It is important to notice that its order of magnitude HjAt is inde- pendent of the amount of the perturbation. The energy change determined by the relation (44.1) will be observed, however weak the interaction between the two parts of the system. This result is peculiar to quantum theory and has a deep physical significance. It shows that, in quantum mechanics, the law a™" The uncertainty relation for energy 151 of conservation of energy can be verified by means of two measurements only to an accuracy of the order of h/At, where At is the time interval between the measurements. The relation (44. 1) is often called the uncertainty relation for energy. How- ever, it must be emphasised that its significance is entirely different from that of the uncertainty relation ApAx ~ h for the co-ordinate and momen- tum. In the latter, Ap and Ax are the uncertainties in the values of the momentum and co-ordinate at the same instant; they show that these two quantities can never have entirely definite values simultaneously. The energies E, e, on the other hand, can be measured to any degree of accuracy at any instant. The quantity (E+e)-(E'+e f ) in (44.1) is the difference between two exactly measured values of the energy E+e at two different instants, and not the uncertainty in the value of the energy at a given instant. If we regard E as the energy of some system and e as that of a "measuring apparatus", we can say that the energy of interaction between them can be taken into account only to within h/At. Let us denote by AE, As, ... the errors in the measurements of the corresponding quantities. In the favour- able case when € , e' are known exactly (Ae = Ae' = 0), we have A(E-E')~hlAt. (44.2) From this relation we can derive important consequences concerning the measurement of momentum. The process of measuring the momentum of a particle (for definiteness, we shall speak of an electron) consists in a collision of the electron with some other ("measuring") particle, whose momenta before and after the collision can be regarded as known exactly.f If we apply to this collision the law of conservation of momentum, we obtain three equa- tions (the three components of a single vector equation) in six unknowns (the components of the momentum of the electron before and after the col- lision). The number of equations can be increased by bringing about a series of further collisions between the electron and "measuring" particles, and applying to each collision the law of conservation of momentum. This) however, increases the number of unknowns also (the momenta of the electron between collisions), and it is easy to see that, whatever the number of col- lisions, the number of unknowns will always be three more than the number of equations. Hence, in order to measure the momentum of the electron, it is necessary to bring in the law of conservation of energy at each collision,' as well as that of momentum. The former, however, can be applied, as we have seen, only to an accuracy of the order of kjAt, where At is the time be- tween the beginning and end of the process in question. To simplify the subsequent discussion, it is convenient to consider an imaginary idealised experiment in which the "measuring particle" is a perfectly reflecting plane mirror; only one momentum component is then of importance, namely that perpendicular to the plane of the mirror. To tailed 1 thC PrCSent analysis h is ° f n0 im POrtance how the energy of the "measuring" particle is ascer- 152 Perturbation Theory §44 determine the momentum P of the particle, the laws of conservation of momentum and energy give the equations p' + P'-p-P = 0, (44.3) \ € ' + E'-e-E\ ~ h/At, (44.4) where P, E are the momentum and energy of the particle, and p, e those of the mirror ; the unprimed and primed quantities refer to the instants before and after the collision respectively. The quantities p, p\ e, e' relating to the "measuring particle" can be regarded as known exactly, i.e. the errors in them are zero. Then we have for the errors in the remaining quantities, from the above equations : AP = AP', AE'-AE ~ hi At. But AE = (BE/dP)AP = vAP, where v is the velocity of the electron (before the collision), and similarly AE' = v'AP' = v'AP. Hence we obtain (v' x -v x )AP x ~hlAt. (44.5) We have here added the suffix x to the velocity and momentum, in order to emphasise that this relation holds for each of their components separately. This is the required relation. It shows that the measurement of the momentum of the electron (with a given degree of accuracy AP) necessarily involves a change in its velocity (i.e. in the momentum itself). This change is the greater, the shorter the duration of the measuring process. The change in velocity can be made arbitrarily small only as At -» oo, but measurements of momentum occupying a long time can be significant only for a free particle. The non-repeatability of a measurement of momentum after short intervals of time, and the "two-faced" nature of measurement in quantum mechanics — the necessity of a distinction between the measured value of a quantity and the value resulting from the process of measurement — are here exhibited with particular clarity, f The conclusion reached at the beginning of this section, which was based on perturbation theory, can also be derived from another standpoint by con- sidering the decay of a system under the action of some perturbation. Let E be some energy level of the system, calculated without any allowance for the possibility of its decay. We denote by t the lifetime of this state of the system, i.e. the reciprocal of the probability of decay per unit time. Then we find by the same method that \E -E-e\~hlr, (44.6) where E, e are the energies of the two parts into which the system decays. The sum £+e, however, gives us an estimate of the energy of the system before it decays. Hence the above relation shows that the energy of a system, t The relation (44.5) and the elucidation of the physical significance of the uncertainty relation for energy are due to N. Bohr (1928). §45 Potential energy as a perturbation 153 in some "quasi-stationary" state, which is free to decay can be determined only to within a quantity of the order of hjr. This quantity is usually called the width Y of the level. Thus T ~ hJT. (447) §45. Potential energy as a perturbation The case where the total potential energy of the particle in an external field can be regarded as a perturbation merits special consideration. The unperturbed Schrodinger's equation is then the equation of free motion of the particle: A0<o) + #ty(o> = 0> k = V(2mElh*) =pjh, (45.1) and has solutions which represent plane waves. The energy spectrum of free motion is continuous, so that we are concerned with an unusual case of perturbation theory in a continuous spectrum. The solution of the problem is here more conveniently obtained directly, without having recourse to general formulae. The equation for the correction 0<*> to the wave function in the first ap- proximation is A^i)+AV (1) = {2mV\W)^\ (45.2) where U is the potential energy. The solution of this equation, as we know from electrodynamics, can be written in the form of retarded potentials, i.e. in the formf ^\x, y, z) = -(m/lnfr) J ^)U{x' t y\ z')e"<r ^v'\r % (45.3) where dV = dx'dy'dz', r 2 = (*-*') 2 +(v-/) 2 +(*-*') 2 . Let us find what conditions must be satisfied by the field Urn order that it may be regarded as a perturbation. The condition of applicability of per- turbation theory is contained in the requirement that 0&> <^ ^<»). Let a be the order of magnitude of the dimensions of the region of space in which the field is noticeably different from zero. We shall first suppose that the energy of the particle is so small that ka is at most of the order of unity. Then the factor e^r i n the integrand of (45.3) is unimportant in an order-of-magnitude estimate, and the integral is of the order of ifjM\U\a 2 , so that iP ~ m\ U\a*i]J®lh 2 , t This is a particular integral of equation (45.2), to which we may add any solution of the same equauon with zero on the right-hand side, i.e. the unperturbed equation (45.1). 154 Perturbation Theory §45 and we have the condition \U\ < h 2 lma 2 (for ka < 1). (45.4) We notice that the expression on the right has a simple physical meaning; it is the order of magnitude of the kinetic energy which the particle would have if enclosed in a volume of linear dimensions a (since, by the uncertainty relation, its momentum would be of the order of hja). Let us consider, in particular, a potential well so shallow that the condition (45.4) holds for it. It is easy to see that in such a well there are no negative energy levels (R. Peierls 1929); this has been shown, for the particular case of a spherically symmetric well, in §33, Problem. For, when E = 0, the unperturbed wave function reduces to a constant, which can be arbitrarily taken as unity: j/» (0) = 1. Since «/» (1) <4 (o) , it is clear that the wave function tf, = 1 +if/<U for motion in the well nowhere vanishes ; the eigenf unction, being without nodes, belongs to the normal state, so that E = remains the least possible value of the energy of the particle. Thus, if the well is suffi- ciently shallow, only an infinite motion of the particle is possible : the particle cannot be "captured" by the well. Attention must be paid to the fact that this result is peculiar to quantum theory; in classical mechanics a particle can execute a finite motion in any potential well. It must be emphasised that all that has been said refers only to a three- dimensional well. In a one- or two-dimensional well (i.e. one in which the field is a function of only one or two co-ordinates), there are always negative energy levels (see the Problems at the end of this section). This is related to the fact that, in the one- and two-dimensional cases, the perturbation theory under consideration is inapplicable for an energy E which is zero (or very small).f For large energies, when ka > 1, the factor e ikr in the integrand plays an important part, and markedly reduces the value of the integral. The solution (45.3) in this case can be transformed; the alternative form, however, is more conveniently derived by returning to equation (45.2). We take as #-axis the direction of the unperturbed motion; the unperturbed wave function then has the form ^ (0) = e ikx (the constant factor is arbitrarily taken as unity). Let us seek a solution of the equation AifW+ktyV = {2m\}i % )TJe ikx in the form ip {1) = e ikx f; in view of the assumed large value of k, it is suffi- cient to retain in A<£ (1) only those terms in which the factor e ikx is differen- tiated one or more times. We then obtain for / the equation 2ik 8/1 dx = 2mUlh 2 , t In the two-dimensional case ^ is expressed (as is known from the theory of the two-dimensional wave equation) as an integral similar to (45.3), in which, instead of e ikr dx'dy'dz'jr, we hwtiirH^Kkr) dx'dy', where H Q W is the Hankel function of the first kind, of zero order, andr 2 = (x— x') + (y—y ) . As k -> 0, the Hankel function, and therefore the whole integral, tend logarithmically to infinity. Similarly, in the one-dimensional case, we have, in the integrand, 2me ikr dx' IK where r = \x— x'\, and as k -> i/*' 1 ) tends to infinity as Ijk. §45 Potential energy as a perturbation 155 whence 0(D = e ikxy = -(i m /^k)e ikx ! Udx. (45.5) An estimation of this integral gives |0&>| ~ mUajfPk, so that the condition of applicability of perturbation theory in this case is \U\ ^ (h 2 jma z )ka = hvja {ka > 1), (45.6) where v = khjm is the velocity of the particle. It is to be observed that this condition is weaker than (45.4). Hence, if the field can be regarded as a perturbation at small energies of the particle, it can always be so regarded at large energies, whereas the converse is not necessarily true.f The applicability of the perturbation theory developed here to a Coulomb field requires special consideration. In a field where U = <x/r, it is impossible to separate a finite region of space outside which U is considerably less than inside it. The required condition can be obtained by writing in (45.6) a variable distance r instead of the parameter a ; this leads to the inequality a//fo <^ 1. (45.7) Thus, for large energies of the particle, a Coulomb field can be regarded as a perturbation.:}: Finally, we shall derive a formula which approximately determines the wave function of a particle whose energy E everywhere considerably exceeds the potential energy U (no other conditions being imposed). In the first approximation, the wave function depends on the co-ordinates in the same way as for free motion (whose direction is taken as the x-axis). Accordingly, let us look for ifs in the form if* = e ikx F, where F is a function of the co- ordinates which varies slowly in comparison with the factor e ikx (but we cannot in general say that it is close to unity). Substituting in Schrodinger's equation, we obtain for F the equation 2ik dF/8x = (2m/# 2 ) UF, (45 . 8) whence if, = e ikx F = constant xe ikx e^/ hv) ^ dx . (45.9) This is the required expression. It should, however, be borne in mind that this formula is not valid at large distances. In equation (45.8) a term AF has been omitted which contains second derivatives of F. The derivative d 2 F/8x 2 , together with the first derivative dFjdx, tends to zero at large distances, but the derivatives with respect to the transverse co-ordinates y and z do not tend to zero, and can be neglected only if x <^ ka 2 . f In the one-dimensional case the condition for perturbation theory to be applicable is given by the inequality (45.6) for all ka. The derivation of the condition (45.4) given above for the three- dimensional case is not valid in the one-dimensional case, owing to the divergence of the resulting function ift^ (see the preceding footnote). % It must be borne in mind that the integral (45.5) with a field U = aijr diverges (logarithmically) when xl\/(y z +z 2 ) is large. Hence the wave function in a Coulomb field, obtained by means of pertur- bation theory, is inapplicable within a narrow cone about the ar-axis. 156 Perturbation Theory §45 PROBLEMS Problem 1. Determine the energy level in a one-dimensional potential well whose depth is small. It is assumed that the condition (45.4) is satisfied. Solution. We make the hypothesis, which will be confirmed by the result, that the energy level \E\ <^ \U\. Then, on the right-hand side of Schrodinger's equation dY/d* 8 = (2tn]h*)[U{x)-Ey,, we can neglect E in the region of the well, and regard ip as a constant, which without loss of generality can be taken as unity: dV/d* 2 = 2m U/h 2 . We integrate this equation with respect to x between two points ±* x such that a <^ x x <^ l//c, where a is the width of the well and k = \/(2m\E\lh 2 ). Since the integral of U(x) converges, the integration on the right can be extended to the whole range from — oo to + oo : d,-*/™- (1) 1 —00 At large distances from the well, the wave function is of the form <\i = e± KX . Substituting this in (1), we find -2k = (2mjk 2 ) j U dx \E\ = (m\2W) h dx We see that, in accordance with the hypothesis, the energy of the level is a small quantity of a higher order (the second) than the depth of the well. Problem 2. Determine the energy level in a two-dimensional potential well U(r) (where 00 r is the polar co-ordinate in the plane) of small depth; it is assumed that the integral J rU dr converges. ° Solution. Proceeding as in the previous problem, we have in the region of the well the equation 1 d / d«/r\ 2m -— (r— ) =—U. r dr\ dr J h 2 Integrating this with respect to r from to r t (where a <^ r x <^ l//c), we find [d</ri 2m } o At large distances from the well, the equation of free motion in two dimensions is 1 d / di£\ 2m r dr\ dry h 2 and has a solution (vanishing at infinity) ifr = constant x Ko(Kr), where Ko is the Hankel function of the first kind, of zero order and imaginary argument; for small values of the argument, the leading term in Ko(*cr) is proportional to log kt. Bearing this in mind, we equate the logarithmic derivatives of ^ for r ~ a inside the well (the right-hand side of (1)) §45 Potential energy as a perturbation 157 and outside it, obtaining 1 2m r =77- U(r)rdr, a log K.a n z a J log whence h* ( h 2 \E\ -exp( [ jlC/jrdrl-A. ma 2 [ m \_j j ) We see that the energy of the level is exponentially small compared with the depth of this well. CHAPTER VII THE QUASI-CLASSICAL CASE §46. The wave function in the quasi-classical case If the de Broglie wavelengths of particles are small in comparison with the characteristic dimensions which determine the conditions of a given problem, then the properties of the system are close to being classical, j- just as wave optics passes into geometrical optics as the wavelength tends to zero. Let us now investigate more closely the properties of "quasi-classical" systems. To do this, we make in Schrodinger's equation V 2m « the substitution = ««/»)* (46.1) For the function a we obtain the equation y — (V a cr)2- V l —H a a = E- U. (46.2) t— 1 2m a *-* 2m„ a a a a Since the system is supposed almost classical in its properties, we seek a in the form of a series : a = ff +(A/«> 1 +(A/«) 8 a 8 + ... , (46.3) expanded in powers of h. We begin by considering the simplest case, that of one-dimensional motion of a single particle. Equation (46.2) then reduces to a'*l2m-iha"l2m = E-U{x), (46.4) where the prime denotes differentiation with respect to the co-ordinate x. In the first approximation we write a = a and omit from the equation the term containing h : (T ' 2 /2m = E- U(x). f We may point out, in particular, that the states of the discrete spectrum with large values of the quantum number n are quasi-classical. For the number n (the ordinal number of the state) determines the number of nodes of the eigenfunction (see §21). The distance between adjoining nodes, however, is of the same order of magnitude as the de Broglie wavelength. For large n this distance is small, so that the wavelength is small in comparison with the dimensions of the region of motion. 158 §46 The wave junction in the quasi-classical case 159 Hence we find c = ± j \/{Zm[E-U{x)]} dx. The integrand is simply the classical momentum p(x) of the particle, expres- sed as a function of the co-ordinate. Defining the function p{x) with the -f sign in front of the radical, we have a =±jpdx, p = ^/[2m{E-U)], (46.5) as we should expect from the limiting expression (6.1) for the wave function.-)- The approximation made in equation (46.4) is legitimate only if the second term on the left-hand side is small compared with the first, i.e. we must have h\<j"la' 2 \ 4 1 or \d(h/a')ldx\ < 1. In the first approximation we have, according to (46.5), a' =p, so that the condition obtained can be written |d(A/27r)/d*| <^ 1, (46.6) where X(x) = 2irhlp(x) is the de Broglie wavelength of the particle, expressed, as a function of x by means of the classical function p(x). Thus we have obtained a quantitative "quasi-classical" condition: the wavelength of the: particle must vary only slightly over distances of the order of itself. The formulae here derived are not applicable in regions of space where this condi- tion is not satisfied. The condition (46.6) can be written in another form by noticing that dft d ntdU m\F\ dx dx p dx p where F = —dUjdx is the classical force acting on the particle in the external field. In terms of this force we find mh\F\/p* <4 1. (46.7) It is seen from this that the quasi-classical approximation becomes inapplic- able if the momentum of the particle is too small. In particular, it is clearly inapplicable near turning points, i.e. near points where the particle, according to classical mechanics, would stop and begin to move in the opposite direction. These points are given by the equation p(x) = 0, i.e. E = U(x). As p -> 0, the de Broglie wavelength tends to infinity, and hence cannot possibly be supposed small. t As is well known, J p dx is the time-independent part of the action. The total mechanical action 5 of a particle is S = —Et± J p dx. The term — Et is absent from o , since we are considering a time- independent wave function tfi. 160 The Quasi-Classical Case §46 Let us now calculate the next term in the expansion (46.3). The first-order terms in % in equation (46.4) give whence a/ = -a "/2a ' = -p'\lp. Integrating, we find o 1 = -ilogp, (46.8) omitting the constant of integration. Substituting this expression in (46.1) and (46.3), we find the wave function in the form tjj = c 1 p- 1 ^ i ^JP dx +C 2 p- 1 ^e^ i / n) JP dx . (46.9) The subsequent terms in the expansion (46.3) lead to the appearance, in the coefficients of the exponentials, of terms in the first and higher powers of h\ it is not usually necessary to calculate these terms. The presence of the factor \jy/p in the wave function has a simple inter- pretation. The probability of finding the particle at a point with co-ordinate between x and x+dx is given by the square |«/f| 2 , i.e. is essentially propor- tional to ljp. This is exactly what we should expect for a "quasi-classical" particle, since, in classical motion, the time spent by a particle in the segment dx is inversely proportional to the velocity (or momentum) of the particle. In the "classically inaccessible" parts of space, where E < U(x), the func- tion p(x) is purely imaginary, so that the exponents are real. The wave func- tion in these regions can be written in the form C ' C ' V\P\ V\P\ ' ' } PROBLEM Determine the wave function in the quasi-classical approximation up to terms of the order of H in the coefficient of the exponent. Solution. The terms of order h 2 in equation (46.4) give <V<V+ W 2 + W = o, whence (substituting (46.5) and (46.8) for a and a ± ) G2 ' =p"l4p2-3p'*l8p3. Integrating (by parts in the first term) and introducing the force F — pp'/tn, we obtain o- 2 = lmF/p*+$m* J* {F*lp) dx. §47 Boundary conditions in the quasi-classical case 161 The wave function in this approximation is of the form = £tlKp = «tf/M».+» 1 (l—^[(j a ) constant r , „ r = [\-limhFlp z -\ihmH (F 2 /p 5 ) dx]eWMb d *. VP J §47. Boundary conditions in the quasi-classical case Let x — a be a turning point, so that U(a) = E, and let U > E for all x > a, so that the region to the right of the turning point is classically inaccessible. The wave function must be damped in this region. Sufficiently far from the turning point, it has the form C / 1 I r |\ tb = expf ad* ) for x > a, (47.1) corresponding to the first term in (46.10). To the left of the turning point, the wave function must be represented by a real combination (46.9) of two quasi-classical solutions of Schrodinger's equation: * = ^KiJH + :S exp H.M for x<a - (47 - 2) a a To determine the coefficients in this combination we must follow the variation in the wave function for positive x—a (where (47.1) holds) to negative x — a. In doing so, however, it would be necessary to pass through a region near the turning point where the quasi-classical approximation is invalid, and the exact solution of Schrodinger's equation must be con- sidered.f This can be avoided if we formally regard f asa function of a complex variable x and go from positive to negative x— a along a path which is always sufficiently far from the point a, so that the quasi-classical condition is formally satisfied. Let us first examine the variation of the wave function (47.1) on passing round the point a from right to left along a semicircle of large radius in the upper half-plane of the complex variable x. For clarity, we may rewrite (47.1) in the form X ftx) = iC[2m(U- E)]-U* exp ( - ^ f y/[U(x) - E] dx\ , (47.3 ) t Near the turning point, E—U^.Fo(x—a) (where Fq = [— dU/dx] x =a); the corresponding exact solution of Schrodinger's equation is given by the formulae derived in §24. 162 The Quasi-Classical Case §48 where the function p(x) has been written explicitly. It is evident a priori that, as a result of passing along the path indicated above, the function (47.3) must become the second term in (47.2), since along the whole of the path this term predominates over the first term, which decreases exponentially into the upper half -plane; and in fact, when x varies along this path, the phase of the difference U(x) — E, like that of x— a, increases by it. Consequently the function (47.3) becomes the second term in (47.2) with coefficient C*2 = \Ce~ in,A . Similarly, on passing from right to left along a semicircle in the lower half-plane, the function (47.3) becomes the first term in (47.2) with coefficient C\ = \Ce inl *. Thus the wave function (47.1) for x > a corresponds for x < a to the functionf ib = cosf- p dx+hr ) V'P VU / a a = sinf- p dx+hr), for x < a. (47.4) Vp \*J / X The functions (47.1) and (47.4) are approximate expressions to the right and left of the turning point for the same exact solution of Schrodinger's equa- tion (H. A. Kramers 1926). If the classically accessible region is bounded (at x = a) by an infinitely high "potential wall", the boundary condition for the wave function at x = a is ifj = (see §18). The quasi-classical approximation is then valid up to the wall itself, and the wave function is 1 C ib = sin- I p dx for x < a, hj C Vp HJ x (47.5) a ib = for x > a. §48. Bohr and Sommerfeld's quantisation rule The results which we have obtained enable us to derive the condition which determines the quantum energy levels in the quasi-classical case. To do this we consider a finite one-dimensional motion of a particle in a potential well : t If the region U < E lay to the right of a turning point x = b, (47.4) would be replaced by b c ib = cos VP (i j p *,+*,) = -£ sing j p d* +i „). (47.4a) §48 Bohr and Sommerf eld's quantisation rule 163 the classically accessible region b ^ x ^ a is bounded by two turning points.f The boundary condition at x = b gives (in the region right of this point) the wave function (47.4a) : = — sirX- J p dx+fr\. (48.1) b Applying formula (47.4) to the region left of the point x = a, we obtain the same function in the form ^ = S sin G/' d * +i 4 VP X If these two expressions are the same throughout the region, the sum of their phases (which is a constant) must be an integral multiple of it: a hi pdX+ilT =(»+1)tT, withC = (-l)"C". Hence ipdx = 2irh(n+i), (48.2) a where §pdx=2$pdxis the integral taken over the whole period of the quasi-classical motion of the particle. This is the condition which determines the stationary states of the particle in the quasi-classical case. It corresponds to Bohr and Sommerfeld's quantisation rule in the old quantum theory. It is easy to see that the integer n is equal to the number of zeros of the wave function, and hence it is the ordinal number of the stationary state. For the phase of the wave function (48.1) increases from ^n at x — b to (« + f)7r at x = a y so that the sine vanishes n times in this range (outside the range b ^ x < a, the wave function decreases monotonically and has no zeros at a finite distance) .% We recall, incidentally, that the quasi-classical f In classical mechanics, a particle in such a field would execute a periodic motion with period (time taken in moving from * = b to x = a and back) a u T = 2 f dxjv = 1m J dxjp, b b where v is the velocity of the particle. J Strictly speaking, the zeros should be counted by means of the exact form of the wave function near the turning points. If this is done, the result given in the text is confirmed. 164 The Quasi-Classical Case §48 approximation, and therefore the quantisation rule (48.2), are applicable only when n is large, f In normalising these wave functions, the integration of J^fJ 2 can be restricted to the range b ^ x ^ a, since outside this range «/r decreases exponentially. Since the argument of the sine in (48.1) is a rapidly varying function, we can with sufficient accuracy replace the squared sine by its mean value \. This gives b rK ' ~ 7rC 2 l2m<o = 1, where w = InJT is the frequency of the classical periodic motion. Thus the normalised quasi-classical function is * = f^ sin [ll pdx+iir } (483) b It must be recalled that the frequency co is in general different for different levels, being a function of energy. The relation (48.2) can also be interpreted in another manner. The integral j> p dx is the area enclosed by the closed classical phase trajectory of the particle (i.e. the curve in the px-plane, which is the phase space of the particle). Dividing this area into cells, each of area 2ttH, we have n cells altogether ; n, however, is the number of states with energies not exceeding the given value (corresponding to the phase trajectory considered). Thus we can say that, in the quasi-classical case, there corresponds to each quantum state a cell in phase space of area 2mh. In other words, the number of states belonging to the volume element ApAx of phase space is ApAx/27rk. (48.4) If we introduce, instead of the momentum, the wave number k =p/h, this number can be written AIiAx/2tt. It is, as we should expect, the same as the familiar expression for the number of proper vibrations of a wave field.J Starting from the quantisation rule (48.2), we can ascertain the general nature of the distribution of levels in the energy spectrum. Let AE be the f In some cases the exact expression for the energy levels E(n) (as a function of the quantum number «), obtained from the exact Schrodinger's equation, is such that it retains its form as n — ► co ; examples are the energy levels in a Coulomb field, and those of a harmonic oscillator. In these cases, of course, Bohr's quantisation rule, although really applicable only for large n, gives for the function E(n) an expression which is the exact one. J See, for example, The Classical Theory of Fields, §52. §48 Bohr and Sommerfeld's quantisation rule 165 distance between two neighbouring levels, i.e. levels whose quantum numbers n differ by unity. Since AE is small (for large n) compared with the energy itself of the levels, we can write, from (48.2), AE i (dp/dE) dx = 2-nh. But dEjdp = v, so that i (dp/dE) dx = i dxjv = T. Hence we have AE = IttH/T = hco. (48.5) Thus the distance between two neighbouring levels is hco. The frequencies co may be regarded as approximately the same for several adjacent levels (the difference in whose numbers n is small compared with n itself). Hence we reach the conclusion that, in any small range of a quasi-classical part of the spectrum, the levels are equidistant, at intervals of hco. This result could have been foreseen, since, in the quasi-classical case, the frequencies cor- responding to transitions between different energy levels must be integral multiples of the classical frequency co. It is of interest to investigate what the matrix elements of any physical quantity / become in the limit of classical mechanics. To do this, we start from the fact that the mean value / in any quantum state must become, in the limit, simply the classical value of the quantity, provided that the state itself gives, in the limit, a motion of the particle in a definite path. A wave packet (see §6) corresponds to such a state ; it is obtained by superposition of a number of stationary states with nearly the same energy. The wave func- tion of such a state is of the form n where the coefficients a n are noticeably different from zero only in some range Aw of values of the quantum number n such that 1 < An <^ n\ the numbers n are supposed large, because the stationary states are quasi-classical. The mean value of/ is, by definition, / = j T*/T d* = 22 a m *a n f mn #»™\ or, replacing the summation over n and m by a summation over n and the difference m —n = s, where we have put co mn = sco in accordance with (48.5). 166 The Quasi-Classical Case §48 The matrix elements f nm calculated by means of the quasi-classical wave functions decrease rapidly in magnitude as the difference m—n increases, though at the same time they vary only slowly with n itself (m—n being fixed). Hence we can write approximately / = SS a*a n f/<»« = 2 K|* S//*—, where we have introduced the notation f s =f- +s -, n being some mean value of the quantum number in the range An. But 2 \a n \ 2 = 1 ; hence dost The sum obtained is in the form of an ordinary Fourier series. Since / must, in the limit, coincide with the classical quantity f(t), we arrive at the result that the matrix elements f mn in the limit become the components f m _ n in the expansion of the classical function f(t) as a Fourier series. Similarly, the matrix elements for transitions between states of the con- tinuous spectrum become the components in the expansion of/(/) as a Fourier integral. Here the wave functions of the stationary states must be normalised by (l//z) times the delta function of energy. All the above results can be generalised immediately to systems with several degrees of freedom, executing a finite motion for which the problem in classical mechanics allows a complete separation of the variables in the Hamilton-Jacobi method (called a conditionally periodic motionf). After separation of the variables for each degree of freedom, the problem reduces to a one-dimensional problem, and the corresponding quantisation conditions are j>pi dqi = 2Trh(ni + yi), (48.6) where the integral is taken over the period of variation of the generalised co-ordinate qi, and yi is a number of the order of unity which depends on the nature of the boundary conditions for the degree of freedom considered .J In the general case of an arbitrary (not conditionally periodic) motion in several dimensions, there are no quantum numbers nt. The concept of cells in phase space is, however, always valid in the quasi-classical approxima- tion. This is clear from the relationship noted above with the number of proper vibrations of the wave field in a given volume of space. In the general t See Mechanics, §50. X For example, in motion in a centrally symmetric field we have j>p r dr = 2Trh(n r +%), j>p d d6 = 2iTh(l—m+\) > j>p^ d<f> = 2-nhm, where n r = n— I— 1 is the radial quantum number. The last of the three equations simply expresses the fact that ptj> is the z-component of the angular momentum, equal to Hm. §49 Quasi-classical motion in a centrally symmetric field 167 case of a system with s degrees of freedom, there are AIV = A 2 i ... A? 5 A£i ... ^p s |(2^^hy (48.7) quantum states in a volume element in phase space. PROBLEMS Problem 1 . Determine (approximately) the number of discrete energy levels of a particle moving in an arbitrary (not central) field C7(r) which satisfies the quasi-classical condition. Solution. The number of states belonging to a volume of phase space which corresponds to momenta in the range ^ p < £max and particle co-ordinates in the volume element dV is f Trpmax 3 dVI(2irh) 3 . For given r the particle can have (in its classical motion) a momentum satisfying the condition E = p 2 /2m + U(r) «£ 0. Substituting />max = V[ — 2m£7(r)], we obtain the total number of states of the discrete spectrum: V (-Ufl*&V, 3tt2 #* where the integration is over the region of space in which U < 0. This integral diverges (i.e. the number of states is infinite) if U decreases at infinity as r~ s with s< 2, in accordance with the results of §18. Problem 2. The same as Problem 1, but for a quasi-classical centrally symmetric field U(r) (V. L. Pokrovskii). Solution. In a centrally symmetric field the number of states is not the same as the number of energy levels, on account of the degeneracy of the latter with respect to the direction of the angular momentum. The required number can be found by noting that the number of levels with a given value of the angular momentum M is the same as the number of (non-degenerate) levels for a one-dimensional motion in a field with potential energy Uett = U(r)+M 2 l2tnr' 2 . The maximum possible value of the momentum p r for given r and energies E =s£ is /v.max = V(— 2mU e ti). The number of states (i.e. the required number of levels) is therefore f^#r = V(2m)r /(_ u _^!\ dr J Ink 2tt% J V V 2mry The required total number of discrete levels is obtained from this by integration with respect to M/H (which replaces in the quasi-classical case the summation with respect to I), and is (m/4£2) j(-U)rdr. §49. Quasi-classical motion in a centrally symmetric field In motion in a centrally symmetric field the wave function of a particle falls, as we know, into an angular and a radial part. Let us first consider the former. The dependence of the angular wave function on the angle <f> (determined by the quantum number m) is so simple that the question of finding approxi- mate formulae for it does not arise. The dependence on the polar angle 9 is, according to the general rule, quasi-classical if the corresponding quantum number / is large (this condition will be more precisely formulated below). 168 The Quasi-Classical Case §49 We shall here confine ourselves to deriving the quasi-classical expression for the angular function for the case (the most important one in applications) of states whose magnetic quantum number is zero (m = 0). This function is, apart from a constant factor, the Legendre polynomial P, (cos0) (see (28.7)), and satisfies the differential equation d 2 Pj/d0 2 +cot d dP t ldd+l(l+ l)Pt = 0. (49.1) The substitution Pi(cos0) = x (0)/Vsin0 (49.2) reduces this to x"+[('+£) 2 +i cosec 2 0] x = 0, (49.3) which does not contain the first derivative and is similar in appearance to the one-dimensional Schrodinger's equation. In equation (49.3), the part of the de Broglie wavelength is played by A = 2tt [(/+$)«+£ cosec^J-i/a. The requirement that the derivative d(A/27r)/da: is small (the condition (46.6)) gives the inequalities 61 > 1, (77-0)/ > 1, (49.4) which are the conditions that the angular part of the wave function is quasi- classical. For large / these conditions hold for almost all values of 0, exclud- ing only a range of angles very close to or n. When the conditions (49.4) are satisfied, we can neglect the second term in the brackets in (49.3) compared with the first: x"+{i+lfx = o. The solution of this equation is x = V sin e ^(cos 0) = A sin[(/+i)0+ a], (49.5) where A and a are constants. For angles <^ 1, we can put in equation (49.1) cot 6 ^ 1/0; replacing also 1(1+1) by the approximation (/+|) 2 , we obtain the equation d 2 P, 1 dPj -+ i +(/+^) 2 P I = 0, d0 2 d0 V ' which has as solution the Bessel function of zero order : P,(cos0) =M(l+i)0]> 6<1- (49.6) The constant factor is put equal to unity, since we must have P z = 1 for 0=0. The approximate expression (49.6) for P, is valid for all angles <^ 1. In particular, it can be applied for angles in the range 1// <^ <^ 1, §49 Quasi-classical motion in a centrally symmetric field 169 where it must agree with the expression (49.5), which holds for all dp 1//. For 61 > 1 the Bessel function can be replaced by its asymptotic expression for large values of the argument, and we obtain 2 8in[(Z+l)0+lir] /- N ttI VO (we can neglect \ in the coefficient compared with /). On comparison with (49.5), we find that A = V(2/w/), a = \n. Thus we obtain finally the following expression for P,(cos 0), applicable in the quasi-classical case:f Pl (cos9) S /- U 7! ; fl . (49.7) V ttI ysinfl The normalised wave function © l0 is obtained, according to (28.7), by multiplying by V"\/(l+\) = i l y/l'. \2 sin[(Z4-i)0+ Jtt] V it ysin0 Let us now turn to the radial part of the wave function. It has been shown in §32 that the function x{ r ) = rR(r) satisfies an equation identical with the one-dimensional Schrodinger's equation, with the potential energy h 2 1(1+1) U l (r) = U(r)+-±-^. 2m r 2 Hence we can apply the results obtained in the previous sections, if the potential energy is understood to be the function C/,(r). The case / = is the simplest. The centrifugal energy vanishes and, if the field U(r) satisfies the necessary condition (46.6), the radial wave function will be quasi-classical in all space. For r = we must have x = 0, and hence the quasi-classical function %(r) is determined by formulae (47.5). If / # 0, the centrifugal energy also must satisfy the condition (46.6). In the region of small r, where the centrifugal energy is of the same order as the total energy, the wavelength A = Qmhjp ~ rjl, and the condition (46.6) gives / > 1 . Thus, if / is small, the quasi-classical condition is violated by the centrifugal energy in the region of small r. It is easily seen that we obtain the correct value of the phase of the quasi-classical wave function %[r) by calculating it from the formulae for one-dimensional motion, replacing the coefficient /(/+1) in the potential energy U^r) by (/+|) 2 4 Ui(r)=U(r)+—±-^-. (49.9) 2m r 2 t Attention is drawn to the fact that, as a result of replacing /(/+ 1) by (/+i) 2 » we have obtained an expression which is multiplied by (—1)' when 6 is replaced by it— 0; this is as it should be for the function Pi(cos 6). % For example, in the simple case of free motion (U = 0) the phase of the function calculated from formula (47.4) with Ui from (49.9) will be kr— JwZ for large r, as it should be. 170 The Quasi-Classical Case §49 The question of the applicability of the quasi-classical approximation to a Coulomb field U = ±a/r requires special consideration. The most import- ant part of the whole region of the motion is that corresponding to distances r for which \U\ ~ \E\, i.e. r ~ cc/|£|. The condition for quasi-classical motion in this region amounts to the requirement that the wavelength A ~ h/ y/(2m\E\) is small compared with the dimensions <x.j\E\ of the region; this gives \E\ ^ ma 2 //* 2 , (49.10) i.e. the absolute value of the energy must be small compared with the energy of the particle in the first Bohr orbit. This condition can also be written in the form d/Hv > 1, (49.11) where v ~ <s/(\E\jm) is the velocity of the particle. It should be noticed that this condition is the opposite of the condition (45.7) for the applicability of perturbation theory to a Coulomb field. The region of small distances ( | U{r) \ > E) is without interest in a repulsive Coulomb field, since for U > E the quasi-classical wave functions diminish exponentially. In an attractive field, however, when / is small it is possible for the particle to penetrate into the region where \U\ p E, so that we have to consider the limits of applicability of the quasi-classical approximation in this case. We use the general condition (46.7), putting there F = -dU/dr = - a/r 2 , p ~ y/(2m\ U\) ~ y/(mijr). As a result, we find that the region of applicability of the quasi-classical approximation is restricted to distances such that r>/* 2 /ma, (49.12) i.e. distances large in comparison with the "radius" of the first Bohr orbit. PROBLEM Determine the behaviour of the wave function near the origin, if the field becomes infinite as ± oc/r s , with s > 2, when r -» 0. Solution. For sufficiently small r, the wavelength A ~ hl\/(m\U\) ~ Hr*l 2 j\/(maL), so that dA/dr ~ hr*' 2 - 1 / V0»°0 <€ 1 1 thus the quasi-classical condition is satisfied. In an attractive field Ui -> — oo when r -> 0. The region near the origin is in this case classically accessible, and the radial wave function x ~ 1/Vp, whence if, ~ y */4-l # In a repulsive field, the region of small r is classically inaccessible. In this case the wave function tends exponentially to zero as r -> 0. Omitting the coefficient of the exponential function, we have T ■HH]' 2 v / (2*«a) exp| — -| | p dr\ |,or ^r~exp{ r -{s/2-i) (s—2)h §50 Penetration through a potential barrier 171 §50. Penetration through a potential barrier Let us consider the motion of a particle in a field of the type shown in Fig. 13, characterised by the presence of a potential barrier, i.e. a region in which the potential energy U(x) exceeds the total energy E of the particle. Fig. 13 In classical mechanics, a potential barrier is "impenetrable" to a particle; in quantum mechanics, however, a particle can pass "through the barrier": the probability of this is not zero (see also §25, Problem 2). If the field U(x) satisfies the quasi-classical conditions, the transmission coefficient for the bar- rier can be calculated in a general form. We may remark that, in particular, these conditions give the result that the barrier must be "wide", and hence the transmission coefficient is small in the quasi-classical case. In order not to interrupt the subsequent calculations, we shall first solve the following problem. Let the quasi-classical wave function in the region to the right of the turning point x = b (where U(x) < E) have the form of a travelling wave : C ri } 1 ip = expl - I p 6.x— \vn I. y/p LHJ J (50.1) We require to find the wave function of this state in the regionf x < b. We shall seek it in the form C" rl) f n ib = exp - p dx\ , V V\P\ Ul J U (50.2) which increases as we go into the region x < b. The exponentially decreasing term is neglected in comparison with the increasing one. To determine the coefficient C we proceed as follows. We notice that, according to formulae | In the problem of penetration through a potential barrier, we are concerned with a motion infinite in both directions; the corresponding levels are doubly degenerate (see §21), and hence the wave functions need not be real. 172 The Quasi-Classical Case §50 (47.1) and (47.4a), there is a correspondence between the functions X 1 ^ = 2VpL eXP |^ J P d *-i £7r )+ ex p{-^ J P <**+!**}] for x > b, (50.3) On the other hand, between two different exact solutions X and 2 of the one- dimensional Schrodinger's equation we have the relation (21.2) tits'— 020i' = constant. We apply this relation with *fj x the solution given by formulae (50.1), (50.2), and 2 the solution (50.3). To the left of the point x = b, we have fcfc'-fck' = -ffl&IW = C7A, while to the right we have Mt-Mi = &W0i)' = -icjh. Equating these two expressions, we obtain C" = — iC. Thus the required quasi-classical wave function is of the form iC for * < b, 1I1 = exp VlPl fl/Hl- c ■■- (50 - 4) for * > 6, = expl- I p dx—liiA. b Let us now go on to calculate the coefficient for the penetration of the barrier by a particle. Let the particle be incident on the barrier from left to right. Since the probability of penetrating the barrier is small in the quasi- classical case, we can with sufficient accuracy write the wave function in region I (Fig. 13), in front of the barrier, the same as if the barrier were completely impenetrable, i.e. in the form (47.4): xf> = — cosl- \p d^+^rj, (50.5) §50 Penetration through a potential barrier 173 where we have introduced the velocity v = p\m\ see below regarding the choice of the normalisation coefficient. If this is written as the sum of two complex expressions, a a the first term (which becomes a plane wave ifi ~ e {i/n)px as x -> -oo) represents a particle incident on the barrier, and the second a particle reflected from the barrier. The normalisation chosen corresponds to a unit probability current density in the incident wave. On the other side of the turning point x = a (in region II, inside the barrier), the wave function (50.5) corresponds, according to the results of §47, to the function x ♦-^[-SHl (50 - 6) a Writing this in the form 6 x i r ii r i ii f n «£ = exp -- pM+t\ PM , (50.7) V VN L h\ J I h\ J U a J> and applying formula (50.4), we find the wave function in region III: b x if, = exp| — I | pdxW- p&x+liir . (50.8) \A> L W J \ nj J v a b The current density in region III, calculated by means of this function, is 6 D=exp [~il \ pdx W' (50 ' 9) Since the current density in the wave incident on the barrier is taken as unity, D is in fact the required transmission coefficient for the barrier. We emphasise that this formula is applicable only when the exponent is large, so that D itself is small. It has been assumed in the foregoing that the field U(x) satisfies the quasi- classical condition over the whole extent of the barrier (excluding only the immediate neighbourhood of the turning points). In practice, however, we often have to deal with barriers where the potential energy curve on one side drops so steeply that the quasi-classical approximation is inapplicable. The 174 The Quasi-Classical Case §50 exponential factor in D remains the same in this case as in formula (50.9), but the coefficient of the exponential (equal to unity in (50.9)) is different. To calculate it we must, essentially, calculate the exact wave function in the "non-quasi-classical" region and determine the quasi-classical wave function inside the barrier in accordance with this. In formula (50.6) a coefficient P # 1 then appears, and in (50.9) a coefficient £ 2 . PROBLEMS Problem 1. Determine the transmission coefficient for the potential barrier shown in Fig. 14: U(x) = for * < 0, U(x) = U —Fx for x > 0; only the exponential factor need be calculated. UM Fig. 14 Solution. A simple calculation gives the result V(2m) D >exp ZhF {U -E)*/j. Problem 2. Determine the probability that a particle (with zero angular momentum) will emerge from a centrally symmetric potential well with U(r) = — U a for r < r„ U(r) = air for r > r (Fig. 15). W Fig. 15 Solution. The centrally symmetric problem reduces to a one-dimensional one, so that the formulae obtained above can be applied directly. We have a/E ■H-ii 7Kr*)] d j §50 Penetration through a potential barrier Evaluating the integral, we finally obtain 175 W r*-> M-m-fi-M'-m In the limiting case r -> 0, this formula becomes W ~ e -{™/nW&m/E) _ e -2ira/nv m These formulae are applicable when the exponent is large, i.e. when ajhv > 1. This condi- tion agrees, as it should, with the condition (49.11) for quasi-classical motion in a Coulomb field. Problem 3. The field U(x) consists of two symmetrical potential wells (I and II in Fig. 16), separated by a barrier. If the barrier were impenetrable to a particle, there would be energy levels corresponding to the motion of the particle in one or other well, the same for both wells. The fact that a passage through the barrier is possible results in a splitting of each of these levels into two neighbouring ones, corresponding to states in which the particle moves simultaneously in both wells. Determine the magnitude of the splitting (the field U(x) is supposed quasi-classical). Solution. Let 2?,, be some level for the motion of the particle in one well (I, say), and ifi (x) the corresponding wave function (so normalised that the integral of ip * over well I is unity). When the small probability of penetration through the barrier is taken into account, the level splits into levels E t and E t with wave functions which are symmetric and anti- symmetric combinations of ^ o 0*0 and 'I'oi—x): M*) = ( W2)[«Ao(*)+0o(-*)]» M*) = (WQM*)-M-*)]- (!) The quasi-classical function tfi (x) diminishes exponentially outside the well, and in particular VW Fig. 16 in the direction of negative x. Hence, within well I, t/> ( _ x) is vanishingly small in compari- son with <Ao(*)> and vice versa in well II. The functions (1) are so normalised that their squares integrated over wells I and II are unity. Schrodinger's equations are ,A "+(2m/£ 2 )(£o- Wo = 0, ^'+(2tnj^){E 1 -U)4, 1 = 0; we multiply the former by <l>i and the latter by ^ , subtract corresponding terms, and integrate 176 The Quasi-Classical Case §50 over * from to oo. Bearing in mind that, for * = 0, ^ = \Z2^ and <Ai' = 0, and that OO 00 J 0o0i <** S — j 0o 2 dx = 1/ V2, we find £i-#o = -(^/^)0o(O)<Ao'(O). Similarly, we find for E 2 —E the same expression with the sign changed. Thus E 2 -E x = (2/z»0 o (O)0 o '(O). By means of formula (47.1), with the coefficient C from (48.3), we find that a ya> r 1 /* n mz;,, — exp|_--j,„d,J, vm—jfm, o where t> = V[2((7 -£ )/m]. Thus a a>/* r 1 r -] E i -E 1 = —exp^-- J |/>| d*J. —a Problem 4. Determine the exact value of the transmission coefficient D for the passage of a particle through a parabolic potential barrier U(x) = — £fe* 2 (supposing that D is mo* small) (E. Kemble 1935).t Solution. Whatever the values of k and E, the motion is quasi-classical at sufficiently large distances |*|, with p = V[2m(E+%kx 2 )] ~ x^(mk)+E^(m/k)fx, and the asymptotic form of the solutions of Schrodinger's equation is = constant xe^ 2 / 2 ^' 6 - 1 / 2 , where we have introduced the notation $ = ximk/h 2 ) 1 ^, e = (EJhW(mlk). We are interested in the solution which, as x -> + oo, contains only a wave which has passed the barrier, i.e. is propagated from left to right. We put as X -> OO, i/j = Be^ 2 / 2 ^- 1 ? 2 , (1) as * -> — 00, ifs = e-i^ a /2|||-ie-i/2 +^£ 8 /2|£|«-i/2 > (2) In the expression (2), the first term represents the incident wave, and the second the reflected wave (the direction of propagation of a wave is that in which its phase increases). The relation between A and B can be found by using the fact that in this case the asymptotic expression for tft is valid in the whole of a sufficiently distant region of the plane of the complex variable £ . Let us follow the variation of the function (1) as we go round a semicircle of large radius in the upper half-plane of £ (cf. §47). Over the whole of this path the term in e*^ 2 is the dominant part of the solution, and hence the function (1) must be converted, by traversing t The solution of this problem can also be applied to penetration sufficiently near the top of any barrier U(x) whose dependence on x near the maximum is quadratic. §51 Calculation of the quasi-classical matrix elements \11 this path, into the second term of the function (2). Hence we find A = B^)"- 1 / 2 = —iBe-™. On the other hand, the condition that the number of particles should be conserved is \A\*+\B\* = 1. From these two relations we find the required transmission coefficient D = \B\*: D = l/(l+e- 2ffe ). This formula holds for any E. When E is large and negative, it gives JD £ e~ mM , in accord- ance with formula (50.9). For E > 0, the quantity R = 1 —D = 1/(1 +e 27te ) is the coefficient of reflection "above the barrier". §51. Calculation of the quasi-classical matrix elements A direct calculation of the matrix elements of any physical quantity/ with respect to the quasi-classical wave functions presents great difficulty. We may suppose that the energies of the states between which the matrix element is calculated are not close to each other, so that the element does not reduce to the Fourier component of the quantity/ ( §48). The difficulties arise because, owing to the fact that the wave functions are exponential (with a large imagin- ary exponent), the integrand oscillates rapidly, and this makes it very troublesome to obtain even an approximate estimate of the integral. We shall consider a one-dimensional case (motion in a field U(x)), and sup- pose for simplicity that the operator of the physical quantity is merely a func- tion/^) of the co-ordinate. Let ^ and «/» 2 be the wave functions correspond- ing to some values E x and E 2 of the energy of the particle (with E 2 > E x , Fig. 17) ; we shall suppose that ^ and if/ 2 are taken real. We have to calculate the integral /l2 = J 0l/& d *« (51.1) \U(x) Fig. 17 178 The Quasi-Classical Case §51 The wave function ifj x in the regions on both sides of the turning point * = a lt but not in its immediate neighbourhood, is of the form (47.1), (47.4a): for x < a lt ifj x C\ -exp[-|j A d,|], W\Pi\ L A| (51.2) for * > fll , if, x = — -L cosf - pj d*— Jfl-V a 1 and similarly for j/t 2 (replacing the suffix 1 by 2). However, the calculation of the integral (51.1) by substituting in it these asymptotic expressions for the wave functions would not give the correct result. The reason is, as we shall see below, that this integral is an exponen- tially small quantity, whereas the integrand is not itself small. Hence even a relatively small change in the integrand will in general change the order of magnitude of the integral. This difficulty can be circumvented as follows. We represent the function ip 2 as a sum if> 2 = ip 2 + +i/if, expressing the cosine (in the region x > a 2 ) as the sum of two exponentials. According to formulae (50.4), we have f or*<* 2) ^ = _=g. exp Q|J fodj( |], x (51.3) the function fa- is the complex conjugate of tfs 2 + : i/t 2 ~ = (iff 2 + )*. The integral (51.1) is also divided into the sum of two complex conjugate integrals /i 2 =/i2 + 4-/i2~, which we shall proceed to calculate. First of all, we note that the integral 00 fl2 + = / 0l/^2 + d* —oo converges. For, although the function «/r 2 + tends exponentially to infinity in the region x < a 2 , the function ifj x , in the region x < a lt tends exponentially to zero still more rapidly (since we have \p x \ > \p 2 \ everywhere in the region x < a 2 ). We shall regard the co-ordinate # as a complex variable, and displace the path of integration off the real axis into the upper half-plane. When x receives a positive imaginary increment, an increasing term appears in the function tp x (in the region * > a^, but the function i/j 2 + decreases still more §51 Calculation of the quasi-classical matrix elements 179 rapidly, since we have p 2 > p x everywhere in the region x > a x . Hence the integrand decreases. The displaced path of integration does not pass through the points * = a lt a 2 on the real axis, near which the quasi-classical approximation is inapplic- able. Hence we can use for i]i x and ip 2 + , over the whole path, the functions which are their asymptotic expressions in the upper half-plane. These are «Ai = expf- [ ^{ImCU-Ei)} &x\ 2[2m(?7-£' 1 )]i/4 F UJ J (51.4) &j + =- — expf — [ ^{2m(U-E 2 )}dx\ where the roots are taken so as to be positive on the real axis for x < a 2 . In the integral — ZCjC2 /ia + = , ,L ? f expl"- fv{2m(C/-£' 1 )} dx— (v{2 m (U-E 2 )} dxl x (51.5) 4V(2m) f(x) dx [(U-Ei)(U-E)]V* we desire to displace the path of integration in such a way that the exponential factor is diminished as much as possible. The exponent has an extreme value only where U(x) = co (for E ± # E 2 , its derivative with respect to x vanishes at no other point). Hence the displacement of the contour of integration into the upper half-plane is restricted only by the necessity of passing round the singular points of the function U(x) ; according to the general theory of linear differential equations, these coincide with the singular points of the wave function ifi(x). The actual choice of the contour depends on the actual form of the field U(x). Thus, if the function U(x) has only one singular point x = x in the upper half-plane, the integration can be effected along the type of path shown in Fig. 18. The immediate neighbourhood of the singular point plays the important part in the integral, so that the matrix element /12 = 2 re/i2 + required is practically proportional to an exponentially small expression of the form /12 - expj- -imf" J ^[2m(E 2 - U)] dx- f V[2m(£i- U)] dal! (51.6) (L. Landau 1932).f t In deriving formulae (51.5) and (51.6), we have replaced the wave functions by their asymptotic expressions, since, in the integral taken along the contour shown in Fig. 18, the order of magnitude of the integral is determined by that of the integrand; hence a relatively small change in the latter does not have any great effect on the value of the integral. 180 The Quasi-Classical Case §51 Fig. 18 The lower limits of the integrals may be any points in the classically accessible regions ; their particular values evidently do not affect the imaginary parts of the integrals. If the function U(x) has several singular points in the upper half-plane, xo in (51.6) must be taken as that for which the exponent is smallest in absolute value.f The quasi-classical matrix elements for motion in a centrally symmetric field must be calculated by the same method. However, we must now replace U(r) by the effective potential energy (the sum of the potential energy and the centrifugal energy), which will be different for states with different /. In view of further applications of the method in question, we shall write the effective potential energies in the two states in a general form, as Ui(r) and Uzir). Then the exponent in the exponential factor in the integrand in (51.5) has an extreme value not only at the points where U\{r) or U2(r) becomes infinite, but also at those where U^-U^r) = E 2 -E v (51.7) Hence, in the formula r r fe~ exd--im[ ( ^[2m(E 2 -U 2 )]dr- JV[2m(£i- U{)] drJJ (51.8) the possible values of r include not only the singular points of U x {r) and C/ 2 ( r )» but also the roots of equation (51.7). The centrally symmetric case differs also in that the integration over r in (51.1) is taken from (and not from -co) to oo: /12 = J Xifx* dr. o Here two cases must be distinguished. If the integrand is an even function t We assume that the quantity f(x) itself has no singular points. §52 The transition probability in the quasi-classical case 181 of r, the integration can be formally extended to the whole range from -oo to oo, so that there is no difference from the previous case. This may occur if U x {r) and U 2 (r) are even functions of r [U(-r) = U(r)]. Then the wave functions Xl (r) and Xi (r) are either even or odd functionsf (see §21), and, if the function /(r) is also even or odd, the product x ±fx% may be even. If, on the other hand, the integrand is not even (as always happens if U(r) is not even), the start of the path of integration cannot be moved away from the point r = 0, and this point must be included among the possible values of r in (51.8). PROBLEMS field*?/ ^Ue-a? 10 " 1 ** thC q uasi - classica l matrix elements (exponential factor only) in a • ^ L YP°£; U(X) becomes infinite °nly for x -> -oo. Accordingly, we put * oo in (51.6). We can extend the integration to + oo. Each of the integrals diverges at the lower limit. Hence we first calculate them from -x to oo, and then pass to the limit x -> oo. We find /l2 r ^ / e^mfaMva-Vi) where vi = V(2fii/«), c* = V(2E 2 lm) are the velocities of the particle at infinity (* -> oo) where the motion is free. " Problem 2 The same as Problem 1, but in a Coulomb field U = et/r, for transitions be- tween states with / = 0. Solution The only singular point of the function U(r) is r = 0. The corresponding integral has been calculated in §50, Problem 2. As a result we have by formula (51.8) f 12 ~ exp — ( — Lh\v 2 vjj §52. The transition probability in the quasi-classical case Penetration through a potential barrier is an example of a process which is entirely impossible in classical mechanics. Another example is "reflection above the barrier" of a particle whose energy exceeds the height of the barrier. In the quasi-classical case the probability of such processes is exponentially small. The relevant exponent can be determined as follows. Considering a transition of any system from one state to another, we solve the corresponding classical equations of motion and find the "path" of the transition; this, however, is complex, in accordance with the fact that the process cannot occur in classical mechanics. In particular, it is found that in general, the "transition point" q at which the formal transition of the system from one state to the other occurs is complex; the position of this point is determined by the classical conservation laws. We next calculate the action Si(?i, q ) + S 2 (q , ? 2 ) for the motion of the system in the first state from some initial position qi to the "transition point" q , and then in the ftl to bS^fa ^rTSLS??^ iS CVen ( ° r ° dd) when ' is even (or odd >' - is seen w ~ 182 The Quasi-Classical Case §52 second state from qo to the final position q<i. The required probability of the process is then given by the formulaf exp im [Si (qi,q ) + S 2 (qo,qz)] • (52.1) If the position of the "transition point" is not unique, it must be chosen so that the exponent in (52.1) has the smallest absolute value (which must yet, of course, be sufficiently large for formula (52.1) to be valid). $ Formula (52.1) is in accordance with the rule derived in §51 for calculating the quasi-classical matrix elements. It should be emphasised, however, that it would not be correct to use the square of the matrix element in calculating the coefficient before the exponential in the probability of such transitions. If formula (52.1) is applied to "reflection above the barrier" of a particle (in the one-dimensional case), qo must be taken as the complex co-ordinate xo of the "turning point" at which the particle reverses its direction of motion, i.e. the complex root of the equation U(x) = E. We shall show how the reflection coefficient can then be calculated more precisely, including the coefficient of the exponential. We must again (as in §50) establish the relation between the wave functions far to the right of the barrier (the transmitted wave) and far to the left (the incident and reflected waves). This is easily done by a method similar to that used in §47, regarding as a function of the complex variable x. We write the transmitted wave in the form X Xi where x± is any point on the real axis, and follow its variation on passing along a path C in the upper half -plane which encloses (at a sufficient distance) the turning point x (Fig. 19) ; the whole of the latter part of this path must lie so far to the left that the error in the approximate (quasi-classical) wave function of the incident wave is less than the required small quantity iff-. Passage round the point x causes a change in the sign of the root \/[E- C/(^)], and after the return to the real axis the function if/+ therefore becomes «/r_, a wave propagated to the left (i.e. the reflected wave). || Since the ampli- tudes of the incident and transmitted waves may be regarded as equal, the t If the transition point is real but lies in the classically inaccessible region, formula (52.1) corres- ponds (in the simple case of one-dimensional motion) to formula (50.9) for the probability of penetra- tion through the potential barrier. An example of the application of formula (52.1) to a problem with several degrees of freedom (the stripping of a deuteron in the field of a nucleus) is given by E. M. Lifshitz, Zhurnal eksperimental'noi i teoreticheskoi fiziki 8, 930, 1938. % If the potential energy of the system has itself singular points, these also must be considered as possible values of go. || A passage along a path below the point *o (simply going along the real axis, for example) converts the function tp+ into the incident wave. §52 The transition probability in the quasi-classical case 183 Fig. 19 required reflection coefficient R is simply the ratio of the squared moduli of «/r_ and ift+ : R = = expf im pdxj. (52.2) Having derived this formula, we can deform the path of integration in the exponent in any manner; if we convert it into the path C shown in Fig. 19, the integral jp dx reduces to twice the integral from x± to x , givingf *0 R = expf im p dxf . (52.3) Xi Since p is real everywhere on the real axis, the choice of x\ does not affect the imaginary part of the integral in the exponent. As already mentioned, among the possible values of xo we must select the one for which the exponent in (52.3) is smallest in absolute magnitude (and this value must be large compared with unity).J It is also implied that, if the potential energy U(x) itself has singularities in the upper half-plane, the integral im- I hj p dx has larger values for such points; otherwise the exponent would be deter- mined by one of these points, but the coefficient of the exponential would not be unity as in (52.3). This condition is certainly not satisfied with increasing t This formula with the coefficient of the exponential (equal to unity) was first obtained by V. L. Pokrovskh, S. K. Sawinykh and F. R. Ulinich (1958). t Of course, only points *o are considered for which x im I p dx > 0, i.e. points lying in the upper half-plane. 184 The Quasi-Classical Case §52 energy E if U{x) becomes infinite anywhere in the upper half-plane: ulti- mately the point xq at which U = E becomes so close to the point #«, where U = oo that the two points give comparable contributions to the reflection coefficient (the integral im- \ p dx ~ 1), and formula (52.3) becomes invalid. In the limit where E is so large that this integral is small compared with unity, perturbation theory becomes applicable (see Problem l).f PROBLEMS Problem 1. Determine the coefficient of reflection above the barrier for particle energies such that perturbation theory is applicable. Solution. Formula (43.1) is used, the initial and final wave functions being plane waves propagated in opposite directions and normalised respectively by unit current density and the delta function of momentum, with dv = dp' and p' the momentum after reflection. Carrying out the integration with respect to p' (taking account of the delta function), we obtain m? I f Wp*\ J (1) —00 This formula is valid if the conditions for perturbation theory to be applicable are satisfied : Ua/Hv <^ 1, where a is the width of the barrier (see the third footnote to §45), and also pa/H £ 1. The latter condition ensures that the function R(p) is not exponential; otherwise the question of the validity of formula (1) would require further investigation. Problem 2. Determine the coefficient of reflection above the barrier for a quasi-classical barrier when the function U(x) has a discontinuity of slope. Solution. If the function U(x) has a singularity for real x, the reflection coefficient is determined mainly by the field near that point, and perturbation theory can be formally applied to calculate it, without having to be valid for all x ; the fulfilment of the quasi-classical condition is sufficient. We then have formula (1) of Problem 1 , the only difference being that the momentum of the incident particle must be replaced by the value ofp(x) at the singular point. In this case we take the point of discontinuous slope as * = 0, and thus have near this point U = — F\x for * > 0, U = —F 2 x for x < 0, with different Fi and F%. The integration with respect to x is effected by including in the integrand a damping factor e ±Xx and then letting A -> 0. The result is m 2 H 2 where po = p(0). t An intermediate case is discussed by V. L. Pokrovskii and I. M. Khalatnikov, Soviet Physics JETP 13, 1207, 1961. §53 Transitions under the action of adiabatic perturbations 185 §53. Transitions under the action of adiabatic perturbations It has already been mentioned in §41 that, in the limit of a perturbation which varies arbitrarily slowly with time, the probability of a transition of a system from one state to another tends to zero. Let us now consider this problem quantitatively, by calculating the transition probability under the action of a slowly varying (adiabatic) perturbation. Let the Hamiltonian of the system be a slowly varying function of time, tending to definite limits as t -> ± oo, and let ifi n (q, t) and E n {t) be the eigen- f unctions and the eigenvalues of the energy (depending on time as a para- meter) obtained by solving Schrodinger's equation i?(*)0» = E n ty n ; on account of the adiabatic variation of i? with time, the time variation of E n and ift n with time will also be slow. The problem is to determine the proba- bility «?i2 of finding the system in a certain state 02 as t -> + oo, if it was in the state 0i as t -> — oo. The slow variation of the perturbation means that the duration of the "transition process" is very long, and therefore the change in the action during this time (given by the integral — J E(t) dt) is large. In this sense the problem is quasi-classical, and the required probability is mainly determined by the values t o of t for which Ei(t ) = E 2 (t ) (53.1) and which correspond, as it were, to the "instant of transition" in classical mechanics (cf. §52); in reality, of course, such a transition is classically impossible, as is shown by the fact that the roots of equation (53.1) are complex. It is therefore necessary to examine the properties of the solutions of Schrodinger's equation for complex values of the parameter t in the neighbourhood of the point t = t o at which the two eigenvalues of the energy become equal. As we shall see, the eigenfunctions 0i, 02 vary rapidly with t near this point. To determine this dependence, we first define linear combinations 01, 02 of 0i, 02 which satisfy the conditions j fa* dq = j 02 2 dq = 0, J* fafa dq = 1. (53.2) This can always be achieved by suitable choice of the complex coefficients (which are functions of t). The functions 0i, 02 have no singularity at t = to. We now seek the eigenfunctions as linear combinations = «i0i+a 2 02. (53.3) Here it must borne in mind that, when the "time" t is complex, the operator fi(t) (of the form (17.4)) is still equal to its transpose (i? = i?), but is no longer Hermitian (i? # i?*), since the potential energy U(t) ^ U*(t). We substitute (53.3) in Schrodinger's equation, multiply on the left by 186 The Quasi-Classical Case §53 fa or fa, and integrate with respect to q. With the notation H tk (t) = jfififadq, (53.4) and using the fact that H±2 = H21 owing to the above-mentioned property of the Hamiltonian, we obtain the equations H n a 1 +H 12 a 2 = Ea 2 , (53.5) Hi 2 a 1 + H 22 a 2 = Ea-y. The condition for these equations to have non-zero solutions is (#12 — E) 2 = H11H22, and the roots of this give the energy eigenvalues E = H 12 ±V(HiiH 22 ). (53.6) Then (53.5) gives a 2 / ai = ±V(HnlH 22 ). (53.7) It is seen from (53.6) that, for a coincidence at the point t = Jo of the two eigenvalues, either Hn or H22 must vanish at that point ; let Hn vanish there. At a regular point, a function in general vanishes as t — to, and therefore E(t) - E(t ) = ± constant x y/(t- 1 ), (53.8) i.e. E(t) has a branch point at t = to. We also have #2 ~ V(* — *o)> an< ^ so there is at the point t = to only one eigenfunction, fa. We now see that the problem is formally completely analogous to the problem of reflection above the barrier discussed in §52. We have a wave function *F(£) which is "quasi-classical" with respect to time, instead of the function quasi-classical with respect to the co-ordinate in §52, and wish to find the term of the form C2fae~ iE £ ln in the wave function for t -> + 00, if the wave function \F(f) = fae~ iE t ln as t -> — 00. This is analogous to the problem of determining the reflected wave for x -> — 00 from the transmitted wave for x -> + 00. The required transition probability w±2 = |c2| 2 - The action S = — J E(t) dt is given by the time integral of a function having complex branch points (just as the function p(x) in the integral j p dx had complex branch points). The problem under consideration is therefore dealt with by means of a contour in the plane of the complex variable t from large negative to large positive values, just as in §52 for the plane of the variable x, and we shall not repeat the derivation here. We shall suppose that E2 > E\ on the real axis. Then the contour must lie in the upper half-plane of the complex variable t (where the ratio g-iEfihje-iEjih increases). The resulting formula (analogous to (52.2)) is «12 = exp ( - - im J E(t) dt\ , (53.9) c where the integration is along the contour shown in Fig. 19 (from left to right). §53 Transitions under the action of adiabatic perturbations 187 On the left-hand branch of this contour E = Ei, and on the right-hand branch E = E 2 . We can therefore write (53.9) in the form to wi2 = exp ( - 2 im j a>2i(t) dtj , (53.10) where o> 2 i = {E^-E^fh, and t± is any point on the real axis of t\ t must be taken as that root of equation (53.1) lying in the upper half-plane for which the exponent in (53.10) is smallest in absolute value.f In addition, besides the direct transition from state 1 to state 2, there may be possible paths through various intermediate states; the probabilities of these are given by analogous formulae. For example, for a transition 1 -> 3 -> 2 the integral in (53.10) is replaced by a sum of integrals: f ( 31 ) t (23) J <*Ki(t)dt+ J oi2s,{t)dt y where the upper limits are the "points of intersection" of the terms #i(*), E 3 (t) and E 3 (t), E 2 (t) respectively. This result is obtained by means of a contour which encloses both these complex points. J t The possible values of t must include points at which E(t) becomes infinite; for such points the coefficient of the exponential in (53.9) will not be unity. t The intermediate states of a continuous spectrum require a special discussion. CHAPTER VIII SPIN §54. Spin Let us consider a system, such as an atomic nucleus, which executes some motion as a whole. We shall suppose that the internal energy of the nucleus has a definite value. The internal state of the nucleus, however, is in general not completely determined by this value; for the "internal" angular mo- mentum L of the nucleus (i.e. the angular momentum of the particles in their motion within the nucleus) may still have various directions in space. The number of different possible orientations of this angular momentum is, as we know, 2L+1. Thus, in considering the motion of the nucleus (in a given internal state) as a whole, we must examine, as well as its co-ordinates, another discrete variable: the projection of its internal angular momentum on some chosen direction in space. Consequently, we see that the formalism of quantum mechanics allows us, in considering the motion of any particle, to introduce, besides its co-ordinates, another variable quantity specific to any given particle, which can take a limited number of discrete values. We have no reason to suppose, a priori, that this variable is absent when the particle is elementary. In other words, we must in general suppose that, in quantum mechanics, some "intrinsic" angular momentum must be ascribed to an elementary particle, regardless of its motion in space. This property of elementary particles is peculiar to quantum theory (it disappears in the limit h -> ; see the second footnote to this section), and hence is essentially incapable of a classical interpretation. In particular, it would be wholly meaningless to imagine the "intrinsic" angular momentum of an elementary particle as being the result of its rotation about "its own axis". The intrinsic angular momentum of a particle is called its spin, as distinct from the angular momentum due to the motion of the particle in space, called the orbital angular momentum.^ The particle concerned may be either elementary, or composite but behaving in some respect as an elementary particle (e.g. an atomic nucleus). The spin of a particle (measured, like the orbital angular momentum, in units of h) will be denoted by s. In the preceding chapters we have always supposed that the three co- ordinates of a particle form a complete set of quantities, so that, if they are given, its state is completely determined. We now see that this is in general t The physical idea that an electron has an intrinsic angular momentum was put forward by G. Uhlenbeck and S. Goudsmit in 1925. Spin was introduced into quantum mechanics in 1927 by W. Pauli. 188 §54 Spin 189 not true : for a complete description of the state of a particle, not only its co- ordinates, but also the direction of the spin vector, must be specified. Hence the wave function of a particle must be a function of four variables : the three co-ordinates, and the spin variable which gives the value of the projection of the spin on a selected direction in space, and takes a limited number of discrete values. These values can be added as a suffix to the wave functions. Thus the wave function of a particle which has a non-zero spin is in fact not one function, but a set of several different functions of the co-ordinates, differing in their spin suffixes. The quantum-mechanical operator corresponding to the spin of the particle, on being applied to the wave function, acts on the spin variable. In other words, it in some way linearly transforms the functions differing only in the spin suffix into one another. The form of this operator will be established later. However, it is easy to see from very general considerations that the operators § x , i y , s z satisfy the same commutation conditions as the operators of the orbital angular momentum. The angular momentum operator is essentially the same as that of an infinitely small rotation. In deriving, in §26, the expression for the orbital angular momentum operator, we considered the result of applying the rotation operator to a function of the co-ordinates. In the case of the spin, this derivation becomes invalid, since the spin operator acts on the spin variable, and not on the co-ordinates. Hence, to obtain the required commutation relations, we must consider the operation of an infinitely small rotation in a general form, as a rotation of the system of co-ordinates. If we successively perform infinitely small rotations about the *-axis and the jy-axis, and then about the same axes in the reverse order, it is easy to see by direct calculation that the difference between the results of these two operations is equivalent to an infinitely small rotation about the #-axis (through an angle equal to the product of the angles of rotation about the x and j-axes). We shall not pause here to carry out these simple calculations, as a result of which we again obtain the usual commutation relations between the operators of the com- ponents of angular momentum ; these must therefore hold for the spin oper- ators also: {$ v J e }=is x , {s z> s x }=is y , {f x ,$ y }=i$z, (54.1) together with all the physical consequences resulting from them. The commutation relations (54.1) enable us to determine the possible values of the absolute magnitude and components of the spin. All the results derived in §27 (formulae (27.7)-(27.9)) were based only on the commutation relations, and hence are fully applicable here also; we need only replace L in these formulae by s. It follows from formula (27.7) that the eigen- values of the sr-component of the spin form a sequence of numbers differing by unity. However, we cannot now assert that these values must be integral, as we could for the component L z of the orbital angular momentum (the derivation given at the beginning of §27 is invalid here, since it was based 190 Spin §54 on the expression (26.14) for the operator l z , which holds only for the orbital angular momentum). Moreover, we find that the sequence of eigenvalues s z is limited above and below by values equal in absolute magnitude and opposite in sign, which we denote by ±s. The difference 2s between the greatest and least values of s e must be an integer or zero. Consequently s can take the values 0, |, 1, f , ... . Thus the eigenvalues of the square of the spin are s 2 = s{s+\), (54.2) where s can be either an integer (including zero) or half an integer. For given s, the component s g of the spin can take the values s, s— 1, ... , — s, i.e. 2s+ 1 values in all. From what was said above, we conclude that the state of a particle whose spin is s must be described by a wave function which is a set of 2s + 1 functions of the co-ordinates, f Experiment shows that the majority of the elementary particles (electrons, positrons, protons, neutrons, jit-mesons and all hyperons (A, 2, H)) have a spin of |. There are also elementary particles, the 7r-mesons and the .K-mesons, whose spin is zero. The total angular momentum of a particle is composed of its orbital angular momentum 1 and its spin s. Their operators act on functions of different variables, and therefore, of course, commute. The eigenvalues of the total angular momentum j = 1+s (54.3) are determined by the same "vector model" rule as the sum of the orbital angular momenta of two different particles (§31). That is, for given values of / and s, the total angular momentum can take the values l+s, l+s— 1, ... , \l— s\. Thus, for an electron (spin |) with non-zero orbital angular momentum /, the total angular momentum can be j = /±|; for / = the angular momentum j has, of course, only the one value/ = \. The operator of the total angular momentum J of a system of particles is equal to the sum of the operators of the angular momentum j of each particle, so that its values are again determined by the vector model rules. The angular momentum J can be put in the form J = L+ S, L = 2 l a , S = 2 s a , (54.4) a where S may be called the total spin and L the total orbital angular momentum of the system. We notice that, if the total spin of the system is half-integral (or integral), the same is true of the total angular momentum, since the orbital angular momentum is always integral. In particular, if the system consists of an even number of similar particles, its total spin is always integral, and therefore so is the total angular momentum. f Since s is fixed for each kind of particle, the spin angular momentum Hs becomes zer o in the limit of classical mechanics (H-> 0). This consideration does not apply to the orbital angular momentum, since I can take any value. The transition to classical mechanics is represented by H ten ding to zero and / simultaneously tending to infinity, in such a way that the product hi remains finite . §55 Spinors 191 The operators of the total angular momentum j of a particle (or J, of a system of particles) satisfy the same commutation rules as the operators of the orbital angular momentum or the spin, since these rules are general com- mutation rules holding for any angular momentum. The formulae (27.13) for the matrix elements of angular momentum, which follow from the com- mutation rules, are also valid for any angular momentum, provided that the matrix elements are defined with respect to the eigenstates of this angular momentum. Formulae (29.7) — (29.10) for the matrix elements of arbitrary vector quantities also remain valid (with appropriate change of notation). PROBLEM A particle with spin J is in a state with a definite value s z = %. Determine the probabilities of the possible values of the component of the spin along an axis z' at an angle 9 to the s'-axis. Solution. The mean spin vector s is evidently along the ar-axis and has magnitude J. Taking the component along the s'-axis, we find that the mean value of the spin in that direction is sT' = i cos 0. We also have s7' = i(w+— w~), where io± are the probabilities of the values Sz- = ±£. Since w++w- — 1, we find w+ = cos 2 J0, w- = sin 2 £0. §55. Spinors Let if/(x, y, z; a) be the wave function of a particle with spin a; a denotes the ^-component of the spin, and takes values from — s to +s. We shall call the functions ip(a) with various values of a the "components" of the wave function. We impose on the choice of these "components" the condition that the integral J |«/r(a) | 2 dV determines the probability that the ^-component of the spin of the particle is equal to a. The probability that the particle is in s an element of volume dV in space is dV 21 |j^(cr) | 2 . If the particle is in a state cr = -s with a definite a-value o- , only the component ifj(o) with a = a is not zero, i.e. the wave function is of the form il*(x,y,z; a) = i/^.jy,*^. In this chapter we shall not be interested in the dependence of the wave func- tion on the co-ordinates. For example, in speaking of the behaviour of the function ifi(a) when the system of co-ordinates is rotated, we can suppose that the particle is at the origin, so that its co-ordinates remain unchanged by such a rotation, and the results obtained will characterise the behaviour of the function ^r(o-) with regard to the spin variable a. Let us effect an infinitely small rotation through an angle 8(f> about the sr-axis. The operator of such a rotation can be expressed in terms of the angular momentum operator (in this case the spin operator), in the form 1 +i8(f> . s s . Hence, as a result of the rotation, the functions ^r(a) become ifj(cr) + Sifj(<r), where 8ip(cr) =i8<f> . s z t[i(a). But f s ^(a) = o^r(cr), so that 8if/(a) = iaift(a)8(f>. By a rotation through a finite angle <f> the functions ift{a) are therefore transformed into ■f(a) =«^(<t). (55.1) 192 Spin §55 In particular, by a rotation through an angle 2tt, they are multiplied by a factor e 27ria , which is the same for all a and is ( - l) 2s (2a is always of the same parity as 2s). Thus we see that, when the system of co-ordinates is com- pletely rotated about an axis, the wave functions of a particle of integral spin return to their original values, while those of a particle of half-integral spin change sign. The variable a differs from the ordinary variables (the co-ordinates) by being discrete. The most general form of a linear operator acting on func- tions of a discrete variable a is (/fl(cr) = S/^aO, (55.2) where the f^ are constants. The parentheses round fift emphasise that the spin argument a which follows does not relate to the original function «/» but to that which results from it when the operator / is applied. It is easy to see that the quantities/^' are the same as the matrix elements of the operator/ defined in the usual manner. For the "eigenfunction" of the operator s z corresponding to the value s z = ct is ip(a) = S^. For this function we have Ao* = g/aA'a,, =/a V The right-hand side of this equation can be rewritten in the form ]}fa f a, 8 aa' and then Aw. = S/a'aA*" (55.3) This equation, however, agrees with the usual definition of the matrix of the operator / with respect to the eigenfunctions of the operator s z . Thus the operators acting on functions of a can be represented in the form of (2*+l)-rowed matrices. In particular, we have for the operators of the spin components themselves (4Mff)= 2(^0(a')» (55.4) and similarlyf for s y > K- According to what has been said above, the matrices s x , s y , s g are identical with the matrices L x , L y , L z obtained in §27, where the letters L and M need only be replaced by s and a. Thus the non- vanishing matrix elements of the spin operators are i s x) a ,a-l = (**)<r-l,a = i\/[(*+ °)(s- (7+ 1)], \ ( s v)o,o-i = -(s y )<r-i, a = -¥V[(s+o)(s-a+l)] t (55.5) In the important case of a spin of \{s = £, a = ± £), these matrices have t Attention is drawn to the fact that the sequence of suffixes in the matrix elements on the right-hand side of equation (55.4) is the reverse of the usual sequence (in (55.3)). §55 Spinors 193 two rows, and are of the form <*•>-€ 3- w "€ "J- w=i C -D- (556) These are called (without the factor ^) Pauli matrices. We may also give the matrices of the complex combinations s ± = s x ±is y : rO In r0 °T (s+) =[o J w =[i 0} (557) By direct multiplication of the Pauli matrices, it is easy to verify that the relations 2s J z = is x , 2s J x = is v , 2s J v = iS t (55.8) hold. Combining these with the usual commutation rules (54.1) we find that sJv+sJ x =0, 4vK4=0, $J g +§J y =Q, (55.9) i.e. the Pauli matrices anticommute with one another. By means of these relations, we can easily verify the following useful formulae : § 2 = |, (55.10) (s.a)(s.b) = i(a.b)+ii§.(axb), (55.11) where a and b are any vectors.f It may be noted that any expression quadratic in the components of s thus reduces to terms independent of s and terms linear in s. Hence it follows that any function of the operator s (of spin \) reduces to a linear function (see Problem 1). Let us consider more closely the "spin" properties of wave functions. When the spin is zero, the wave function has only one component, ^(0). When the spin operators act upon it, the result is zero : S X lp = Sylfj = S z *ji = 0. Since the spin operators are related to the rotation operators, this means that the wave function of a particle with spin zero is invariant under rotation of the co-ordinate system, i.e. it is a scalar. The wave functions of particles with spin \ have two components, i}j{\) and ^( — £). For convenience in later generalisations, we shall call these compo- nents i/j 1 and i/j 2 respectively (with upper indices 1 and 2). In any rotation of the co-ordinate system, tp 1 and ip 2 undergo a linear transformation : 0v = a^+j^ 2 , V 2 ' = rl> x +W- (55.12) t The terms on the right which are independent of s must, of course, be understood as constants multiplying the unit two-by-two matrix. 194 Spin §55 The coefhcientsf a, jS, y, 8 are in general complex and functions of the angles of rotation. They are connected by a relation which we derive by considering the bilinear form 0^-0^1, (55.13) where (ifi 1 , xjj % ) and (<£\ cf> 2 ) are two wave functions transformed according to (55.12). A simple calculation gives 0i'02'-02'0i' = ( a 8- i 8y)(^ 2 -^i) ) i.e. the quantity (55.13) is transformed into itself when the co-ordinate system is rotated. If, however, there is only one function which is transformed into itself, it can be regarded as corresponding to zero spin, and therefore must be a scalar, i.e. must remain unchanged when the co-ordinate system is rotated in any manner. Hence we have aS-j8y = 1. (55.14) This is the required relation. The linear transformations (55.12) which leave the bilinear form (55.13) invariant are called binary transformations. A quantity having two compo- nents which undergoes a binary transformation when the co-ordinate system is rotated is called a spinor. Thus the wave function of a particle with spin \ is a spinor. It is possible to put the algebra of spinors in a form analogous to that of tensor algebra. This is done by introducing a vector space of two dimen- sions, in which the metric is defined by an antisymmetrical "metric tensor": The vectors in this space are spinors. Besides the contravariant components 0\ ^ 2 of the spinor, we may introduce the covariant components in accordance with the usual formulae of tensor algebra : fa = 5«v^, so that fa = <A 2 , fa = -V- (55.16) The binary transformations for the covariant components of a spinor are obviously of the form fa' = Zfa-yfa, fa' = -j3<Ai+a«A 2 . (55.17) t Called the Cayley-Klein parameters. §55 Spinors 195 The converse transformation from covariant to contravariant components can be written in the form «A A = S^ty M , (55.18) where the contravariant "metric tensor" g^ has the components which are the same as the components g^. The invariant combination (55.13) can be written as a "scalar product" +Hx = Wi+tfVi = W-0V = 5VA A ^; (55.20) here, and in what follows, summation is implied over repeated {dummy) indices, as in tensor algebra. We may note the following rule which has to be borne in mind in spinor algebra. We have tf Va = Wi+tfVi = -M 2 -^ 1 = -M A - Thus *Va = ~^ A - (55.21) Hence it is evident that the scalar product of any spinor with itself is zero : ^ = 0. (55.22) The expression |^l|2 + |02|2 = ^l*^^ which gives the probability of finding the particle at a given point in space, must clearly be a scalar. Comparing it with the scalar (55.20), we see that the components j/t 1 *, 2 * of the wave function which is the complex conjugate of j/t 1 , ifs 2 are transformed as covariant components of a spinor, i.e. as ip 2 ,—^ 1 respectively : 0i*' = 80 1 *— yifi 2 *, if> 2 *' = — j80 1# +a^ 2 *. On the other hand, by taking the complex conjugate equations to (55.12) 0i*' = a*^ 1 *-^*^ 2 *, 2 *' = y*0!*+8*0 2 * and comparing them with the above, we find that the coefficients a, jS, y, S are related also by a =5*, j8 = -y*. (55.23) By virtue of the relations (55.14), (55.23), the four complex quantities a, jS, y, S actually contain only three independent real parameters, correspond- ing to the three angles which define a rotation of a three-dimensional system of co-ordinates. 196 Spin §56 The fact that 1 *, iff 2 * are transformed as if/ 2 , —if* 1 is closely related to the symmetry with respect to a change in the sign of the time. As was remarked in §18, in quantum mechanics a change in the sign of the time corresponds to a replacement of the wave function by its complex conjugate. When the sign of the time is changed, however, so is that of the angular momentum. Hence the functions which are the complex conjugates of the components 1 , ifj 2 corresponding to projections of the spin a = \ and a = —\ must be equivalent in their properties to the components corresponding respectively to projections of the spin a = — \ and a = \. PROBLEMS Problem 1. Reduce an arbitrary function of the scalar a+2b . s (where s is the operator of spin i) to another linear function of s. Solution. To determine the coefficients in the required formula f(a + 2b . s) = oc+2/?b . s/b, we note that, when the sr-axis is taken in the direction of b, the eigenvalues of the operator a + 2b . s are a ± b, and the corresponding eigenvalues of the operator f(a +2b . s) are/(a±6). Hence we find a = h[f(a + b)+f(a-b)], j3 = £[/( a +&)_/( a _£)]. Problem 2. Determine the values of the scalar product si . S2 of spins (£) of two particles in states in which the total spin of the system, S = Si +S2, has definite values (0 or 1). Solution. From the general formula (31.2), which is valid for the addition of any two angular momenta, we find si . S2 = i for S = 1, Si . S2 = —J for S = 0. Problem 3. Which powers of the operator s of an arbitrary spin s are independent? Solution. The operator (S z -s)(s z -s+l) ... (s z + s), formed from the differences between S z and all possible eigenvalues s z , gives zero when it is applied to any wave^ function, and is therefore itself zero. Hence it follows that (S z ) 28+1 is expressed in terms of lower powers of the operator S z , so that only its powers from 1 to 2s are independent. §56. Spinors of higher rank Analogously to the transition from vectors to tensors in ordinary tensor algebra, we can introduce the idea of spinors of higher rank. Thus, a quantity ip^ t having four components which are transformed as the products j/t\^* of the components of two spinors of rank one, is called a spinor of rank two. Besides the contravariant components «/rV we can consider the covariant components j/t^ and the mixed components 0^ which are transformed as the products ifix<f>n and ipx^ respectively. The transition from one set of components to another is effected by means of a "metric tensor" gx „, in accordance with the usual formulae Thus ip 12 = —ip! 1 = —iff 21 , 0ii = *Pi 2 = ip 22 , and so on. Spinors of any rank are similarly defined. The quantities g^ themselves form an anti- symmetrical spinor of rank two. It is easy to see that the values of its com- ponents remain unchanged under binary transformations. §56 Spinors of higher rank 197 It is easily verified that the product g\ v g^ v is, as it should be, a unit spinor of rank two, i.e. a spinor with components 8\ = S| = 1, Sf = S2 = 0. Thus ibF* = V- (56.1) As in ordinary tensor algebra, there are two fundamental operations in spinor algebra: multiplication, and contraction with respect to a pair of in- dices. The multiplication of two spinors gives a spinor of higher rank ; thus, from two spinors of ranks two and three, iff^ and ip v P a , we can form a spinor of rank five, i/fXfi l P vpa ' Contraction with respect to a pair of indices (i.e. sum- mation of the components over corresponding values of one covariant and one contravariant index) decreases the rank of a spinor by two. Thus, a contraction of the spinor ^x^ 9 " with respect to the indices ju and v gives the spinor ^A^ po °f rank three; the contraction of the spinor ^y* gives the scalar tp\*. Here there is a rule similar to that expressed by formula (55.21): if we interchange the upper and lower indices with respect to which the contraction is effected, the sign is changed (i.e. ^ A = — if**\). Hence, in particular, it follows that, if a spinor is symmetrical with respect to any two of its indices, the result of a contraction with respect to these indices is zero. Thus, for a symmetrical spinor ijj X/l of rank two, we have ift\^ = 0. A spinor of rank n symmetrical with respect to all its indices is called a symmetrical spinor of rank n. From an asymmetrical spinor we can construct a symmetrical one by the process of symmetrisation, i.e. summation of the compo- nents obtained by all possible interchanges of the indices. From what has been said above, it is impossible to construct (by contraction) a spinor of lower rank from the components of a symmetrical spinor. Only a spinor of rank two can be antisymmetrical with respect to all its indices. For, since each index can take only two values, at least two out of three or more indices must have the same value, and therefore the compo- nents of the spinor are zero identically. Any antisymmetrical spinor of rank two is a scalar multiple of the unit spinor g^. We may notice here the fol- lowing relation : gAv&v+g^X+g^ = (56.2) (where ip\ is any spinor), which follows from the above; this rule is simply a consequence of the fact that the expression on the left is (as we may easily verify) an antisymmetrical spinor of rank three. The spinor which is the product of a spinor j/^ with itself, on contraction with respect to one pair of indices, becomes antisymmetrical with respect to the other pair: Hence, from what was said above, this spinor must be a scalar multiple of the spinor gx^. Defining the scalar factor so that contraction with respect to the second pair of indices gives the correct result, we find hA v = -W*^Tsv ( 56 - 3 ) 198 Spin §57 The components of the spinor ^ M * which is the complex conjugate of f/fXfi... are transformed as the components of the contravariant spinor tp^f 1 —, and conversely. The sum of the squared moduli of the components of any spinor is consequently invariant. §57. The wave functions of particles with arbitrary spin Having developed a formal algebra for spinors of any rank, we can now turn to our immediate problem, to study the properties of wave functions of particles with arbitrary spin. This subject is conveniently approached by considering an assembly of particles with spin |. The greatest possible value of the ^-component of the total spin is |w, which is obtained when s g = \ for every particle (i.e. all the spins are directed the same way, along the #-axis). In this case we can evidently say that the total spin S of the system is also \n. All the components of the wave function i/t(a lt cr 2 , ... , a n ) of the system of particles are then zero, except for «/r(|, \, ... , §). If we write the wave function as a product of n spinors if/ty ... , each of which refers to one of the particles, only the component with A, p, ... = 1 in each spinor is not zero. Thus only the product j/r 1 ^ 1 ... is not zero. The set of all these products, however, is a spinor of rank n which is symmetrical with respect to all its indices. If we transform the co-ordinate system (so that the spins are not directed along the #-axis), we obtain a spinor of rank n, general in form except that it is symmetrical as before. The "spin" properties of wave functions, being essentially their properties with respect to rotations of the co-ordinate system, are evidently identical for a particle with spin 5 and for a system of n = 2s particles each with spin \ directed so that the total spin of the system is s. Hence we conclude that the wave function of a particle with spin s is a symmetrical spinor of rank n = 2s. It is easy to see that the number of independent components of a sym- metrical spinor of rank 2s is equal to 2s +1, as it should be. For all those components are the same whose indices include 2s ones and twos; so are all those with 2s— 1 ones and 1 two, and so on up to ones and 2s twos. Mathematically we can say that the symmetrical spinors provide a classifica- tion of the possible types of transformation of quantities when the co-ordinate system is rotated. If there are 2s +1 different quantities which are transformed linearly into one another (and which cannot be reduced in number by any choice of linear combinations of them), then we can assert that their law of transformation is equivalent to that of the components of a symmetrical spinor of rank 2s."f" Any set of any number of functions which are transformed linearly into one another when the co-ordinate system is rotated can be reduced (by an appropriate linear transformation) to one or more symmetrical spinors. f In other words, the symmetrical spinors form what are called irreducible representations of the rotation group (see §98). §57 The wave functions of particles vdth arbitrary spin 199 Thus an arbitrary spinor ^A/xv... °f rank n can De reduced to symmetrical spinors of ranks », n — 2, n — 4, ... . In practice, such a reduction can be made as follows. By symmetrising the spinor j/^... with respect to all its indices, we form a symmetrical spinor of the same rank n. Next, by contracting the original spinor j/^v... with respect to various pairs of indices, we obtain spinors of rank n— 2, of the form «/f A Av , which, in turn, we symmetrise, so that symmetrical spinors of rank n—2 are obtained. By symmetrising the spinors obtained by contracting j/^... w i tn respect to two pairs of indices, we obtain symmetrical spinors of rank n— 4, and so on. We have still to establish the relation between the components of a sym- metrical spinor of rank 2s and the 2s +1 functions ^(cx), where a — s, s— 1, ... , —s. The component .11. ..1 22. ..2 s+a s—cr in whose indices 1 occurs s+a times and 2 s— a times, corresponds to a value a of the projection of the spin on the sr-axis. For, if we again consider a system of n = 2s particles with spin ^, instead of one particle with spin s, the product tfjty 1 . . . x 2 p 2 - • • corresponds to the above component; this product belongs to a s+a state in which s+a particles have a projection of the spin equal to £, and s — a a projection of — $, so that the total projection is ?(s + a) — %(s — a) = a. Finally, the proportionality coefficient between the above component of the spinor and ifj(o) is chosen so that the equation J s |*(a)|»- jy**-|« ( 57 -!) holds; this sum is a scalar, as it should be, since it determines the probability of finding the particle at a given point in space. In the sum on the right-hand side, the components with (s + a) indices 1 occur (2*)! (,+ ff )!(,-a)l times. Hence it is clear that the relation between the functions ip(a) and the components of the spinor is given by the formula 0( ^ = /r (2 * )! V 1 - 1 22 - 2 (57 2) G) V U+a)!(*-<?)U s+ff *- a ' The relation (57.2) ensures the fulfilment not only of the condition (57.1), but also, as we easily see, of the more general condition «A v ~£w. = 5 (- 1) 8 -^)^- *)> ( 57 - 3 ) where ifj^'" and <f>Xa... are two different spinors of the same rank, while 200 Spin §58 iff(<j),(f>(o) are functions derived from these spinors by formula (57.2); the factor ( — l) s -° is due to the fact that, when all the indices of the spinor com- ponents are raised, the sign changes as many times as there are twos among the indices. Formulae (55.4) determine the result of the action of the spin operator on the wave functions ip(a). It is not difficult to find how these operators act on a wave function written in the form of a spinor of rank 2s. For a spin £, the functions i/f(£), «/r( — |) are the same as the components ^r 1 , 2 of the spinor. According to (55.4) and (55.6), the result of the spin operators' acting on them will be (40) x =# 2 , (W = -^ 2 , (W=#\ (40) 2 = W, {Syrfsf = W, (U) 2 = -W- 7A) To pass to the general case of arbitrary spin, we again consider a system of 2s particles with spin |, and write its wave function as a product of 2s spinors. The spin operator of the system is the sum of the spin operators of each particle, acting only on the corresponding spinor, the result of this action being given by formulae (57.4). Next, returning to arbitrary symmetri- cal spinors, i.e. to the wave functions of a particle with spin s, we obtain , g ..11. ..22... .. N .11... 82... ... . 11 ... 22 s+a 8-a ' s+a-1 s-a+1 ' T 8+a+l s-a-X , g ..11. .22... ... ..11... 22... . ..11... 22 8+a 8-a v ' 8+a-l s-a+1 v ' r 8 +o+l s-a-1 ,a ,.H •••22 ... .11 ...22... (**#) = # • s+o s-a s+a 8-a (57.5) We notice that, by starting from these formulae and the relations (57.2), we could derive the expressions (55.5) for the matrix elements of the spin operator acting on the functions ift(o). §58. The relation between spinors and tensors Hitherto we have spoken of spinors as wave functions of the intrinsic angular momentum of elementary particles. Formally, however, there is no difference between the spin of a single particle and the total angular momentum of any system regarded as a whole, neglecting its internal structure. It is therefore evident that the transformation properties of spinors apply equally to the behaviour, with respect to rotations in space, of the wave functions ipj m of any particle or system of particles with total angular momentum /, independent of whether orbital or spin angular momentum is concerned. There must therefore be some definite relation between the laws of transformation for the eigen- functions ^ m under rotations of the co-ordinate system and those for the components of a symmetrical spinor of rank 2/. §58 The relation between spinors and tensors 201 In establishing this relation we must, however, make a clear distinction between two aspects of the dependence of the wave functions on the component m (for a given value of j). The wave function may be regarded as the probability amplitude for various values of m, or may be considered for a given value of m. These two aspects have already been discussed at the beginning of §55, where we dealt with the "eigenf unction" S^ of the operator s z which corres- ponds to s z = <to. The mathematical difference between them is especially clear for a particle of spin s = |. In this case the spin function is, with respect to the variable a, a contravariant spinor of rank 1, i.e. must be written in spinor notation as 8°"^ . With respect to cto it is therefore a covariant spinor. This is evidently a general result: the eigenfunctions 0y m can be put in correspondence with the components of a covariant symmetrical spinor of rank 2/ by means of formulae analogous to (57.2) :f 0/m = — — -011... 22...- (58.1) V (i + tn)\(i — m)\ The eigenfunctions of integral angular momentum j are spherical harmonics, and formula (58.1) relates them to the components of a covariant spinor of even rank. The case; = 1 is of particular importance. The three spherical harmonics Y lm are J^ee^ = TiJ^n x± in v)t 3 .. /3 Fi ± i = T i ' where n is a unit vector along the radius vector. Comparing with (58.1), we see that the components of a spinor of rank two can be brought into corres- pondence with the components of some vector by the formulae J i i * 012 = a z , 0n= -A a x+ia y ), 022 = — -riflx—fay), ( 58 - 2 ) V 2 v 2 v 2 t This result can also be regarded somewhat differently. If the wave function ^ of a particle in state with angular momentum j is expanded in terms of the eigenfunctions tftj m : ifj = S a m ifij m> m then the coefficients a m are the probability amplitudes for various values of m. In this sense they correspond to the "components" t/)(m) of a spin wave function, and this gives their law of transformation. On the other hand, the value of at a given point in space cannot depend on the choice of the co- ordinate system, i.e. the sum S a m iffj m must be a scalar. Comparing with the scalar (57.3), we see that a m must transform as (— \) i ~ m 'jsj ,-m- % Here the functions ifsj m and the components of the vector are related by 0io = ia z , 0n= —(ax+iay), 01 -1 = — r(%-«fy)- (58.2a) V 2 V 2 202 Spin §58 or V = - -Z&> <£ U = —z{a x -ia y \ <p = _ Jt ax+iay y (58 3) Conversely a z = iV20i2, «* = -i-(022_0ii) 5 % = ( 0ii + 022). (58>4) v^ V 2 It is easily verified that with these definitions we have </v£ A /<= a.b, (58.5) where a and b are vectors corresponding to the symmetrical spinors i/t^ and (f)^. It is also not difficult to see that the spinor ift/<f>t iv + 0/0 A " corresponds to the vector <y/2a x b. The relations (58.2) or (58.3) are a particular case of a general rule: any sym- metrical spinor of even rank 2j, where j is integral, can be correlated with a symmetrical tensor of half the rank (j) which gives zero on contraction with respect to any pair of indices; we call this an irreducible tensor. This follows from the fact that the numbers of independent components of the spinor and of the tensor are the same (2/+ 1), as may easily be seen.f The relation between the components of the spinor and of the tensor can be found by means of formulae (58.2) — (58.4), if we consider a spinor of the rank concerned as the product of several spinors of rank two, and the tensor as a product of vectors. Finally, let us determine the relation between the angles of rotation of the co-ordinate system and the coefficients <x, /?, y, 8 of the binary transformation. This is done by noticing that, on the one hand, the cosines of the angles between the original and final axes of co-ordinates are the coefficients in the formulae for the transformation of the components of a vector: <*'t =g*afi* (58.6) and, on the other hand, this same transformation can be performed by means of a binary transformation, using formulae (58.4), (55.12). Thus, for instance, we have a' z =*V 2 «A 12 ' =^V 2 [ a y^ 1 +W 22 +(a3+ i 8y)0 12 ] = (-ay+)88)a e +i(ay+iS8)fl,+(a8+j8y)fl. We can similarly determine the remaining oi ik , and thus obtain the following scheme of transformation coefficients : («-ik) = l (a 2_£2_ y 2 + S 2) ^(_ a 2_ / J2 + y 2 + S 2) (-a/J+yg)" ^ai-jgi+yi-S") Ka^+yHS 2 ) -/(ajS+yS) (-ay+jSS) *(«y+j88) ocS+Py (58.7) t Mathematically we can say that the 2/+ 1 components of an irreducible tensor of rank j (an integer), the 2/+1 spherical harmonics Yj m , and the 2/+1 components of a symmetrical spinor of rank 2/ give the same irreducible representation of the rotation group. §58 The relation between spinors and tensors 203 The inverse expressions for the coefficients a, jS, y, 8 in terms of the angles of rotation of the co-ordinate system can be found by using the Eulerian angles to define the rotation. The matrix of the coefficients a, /?, y, S for a rota- tion through an angle <f> about the sr-axis (denoted by co(<f>)), according to formula (55.1) with cr = + J, is of the form The matrix Q(0) expressing a rotation through an angle 6 about the x-axis is easily calculated from formulae (58.7), in which oc X x = 1> <*-yy — <*zz — cos #> V-yz = ~ a zy = sin "> 'X-xy = Kyx = a #z = &-zx = : [cosA0 *sinA0~| • • i/i I ' ism*v cos*0J (58.9) A rotation specified by the Eulerian angles <f>, 6, ifs (Fig. 20 ; OAT is the line of intersection of the xy and x'y' planes) is carried out in three stages : a rotation *>K through an angle <f> about the #-axis, one through an angle 6 about the new position of the x-axis, and finally one through an angle tfj about the final direc- tion of the #-axis. Accordingly, the matrix of the complete transformation is equal to the product co(i[t)Cl(d)co(<f>). By direct multiplication of the matrices, we finally obtain Ly 8J \j cos^.e^ + *)/ 2 sinA0.e^-* )/2 i sin \Q COSifl g-£(0-0)/2' g -t(^+0)/2 ] (58.10) 204 Spin §59 In particular, a rotation through an angle it about the jy-axis corresponds to Eulerian angles = 7r, 0-0 = tt, so that a = S = 0, j8 = 1, y = — 1. This means that the spinor components are transformed by such a rotation according to0!' = 08,02' = -01, or 0i' = 0!, 02' = 2 . (58.il) PROBLEMS Problem 1. Rewrite the definition (57.4) of the operator of spin J in terms of the spinor components of the vector s. Solution. By means of formulae (58.3), which give the relation between the vector s and the spinor S X|X , the definition (57.4) can be written as i 2\/2 Problem 2. Derive formulae which determine the effect of the spin operator on a vector wave function of a particle with spin 1 . Solution. The relation between the components of the vector function l|> and the com- ponents of the spinor ^ is given by formulae (58.3), and from (57.5) we have 40+ — -0+, S z tfi- = 0_, s z i/j z = or 403 = -*"02/, S z lf> v = tlfl z , 402 = 0. The remaining formulae are derived from these by cyclic permutation of the suffixes x, y, z. They can be written together as 40* = —iemvph The complex vector %\> can be put in the form l|> = e ia (u+iv), where u and v are real vectors, which can be taken to be mutually perpendicular if the common phase a is suitably chosen. The two vectors u and v determine a plane which has the property that the spin component perpendicular to it can take only the values ± 1 . §59. Partial polarisation of particles By a suitable choice of the direction of the #-axis, we can always cause one component (e.g. 2 ) of a given spinor A , the wave function of a particle with spin ^, to vanish. This is evident from the fact that a direction in space is determined by two quantities (angles), i.e. the number of disposable parameters is just equal to the number of quantities (the real and imaginary parts of the complex 2 ) which it is desired to make zero. Physically this means that, if a particle with spin \ (for definiteness, we shall speak of an electron) is in a state described by a spin wave function, then there is a direction in space in which the component of the particle spin has the definite value a = \. We can say that in such a state the electron is completely polarised. There are also, however, states of an electron which may be said to be partially polarised. Such states are not described by wave functions but only by density matrices, i.e. they are mixed states (with respect to spin) (see §14). §59 Partial polarisation of particles 205 The spin density matrix of an electron is a spinor p A <" of rank two normalised by the condition p/ = P1 i+ P2 2 = 1, (59.1) and satisfying the "Hermitian" condition (pa")* = p/- (59.2) For a pure (i.e. completely polarised) spin state of the electron the spinor p A -" reduces to a product of components of the wave function «/r A : Px p = (0A)#^. (59.3) The "diagonal" components pi 1 and p2 2 of the density matrix determine the probabilities of the values + 1 and — f of thcsr-component of the electron spin. The mean value of this component is therefore ^ = KP1 1 — /»2 2 )i or, using (59.1), Pi 1 = H*> P2 2 = i-Tz- (59.4) In a pure state the mean value of the quantities s ± = s x ± is y is calculated as sL = ifjtes-ifj*- = ift 2 *ifj l (see (55.7)). Accordingly we have in a mixed state Pi 2 = n, P2 1 = r-. (59.5) Thus we see that all the components of the spin density matrix of the electron are expressed in terms of the mean values of components of its spin vector. In other words, the real vector s entirely determines the polarisation properties of a particle with spin J. In the limit of complete polarisation one of the com- ponents of this vector (with an appropriate choice of the directions of the axes) is ^ and the other two are zero. In the opposite case of an unpolarised state all three components are zero. In the general case of an arbitrary partial polarisation and any choice of the co-ordinate system we have ^ p ^ 1, where p = 2(^2 + ^2 + sl 2)l/2 is a quantity which may be called the degree of polarisation of the electron. For a particle of arbitrary spin s, the density matrix is a spinor p\ /l ... f " T -" of rank 4s, symmetrical in the first 2s and the last 2s indices and satisfying the conditions Pa,../*- = 1, (59.6) (PA/*...^-)* = P^...^-. (59.7) To calculate the number of independent components of the density matrix, we note that, among the possible sets of values of the indices A, /*, ... (or />, a, . . .) 206 Spin §60 there are only 2s +1 which are essentially different. Using also the fact that the components of the spinor p X/l ... p(T — are related by (59.6), we find that the number of different components is (2s+l) 2 -l = 4s(s+l). Although these components are complex, the relation (59.7) shows that this does not increase the total number of independent quantities describing the state of partial polarisation of the particle, which is therefore 4s(s+l).f For comparison, it may be remarked that the state of complete polarisation of the particle is de- scribed by only 4s quantities (the 2s + 1 complex components of the wave function j/r'M"-, related by one normalisation condition and containing one common phase which is unimportant in the description of the state). Like any spinor of rank 4s, the spinor ^..Z '— is equivalent to a set of irreducible tensors of ranks 4s, 4s -2, ... , 0. In the present case there is only one tensor of each rank, since, on account of the symmetry properties of the spinor ^Xfi... p(T "'i eacn contraction of it can be carried out in only one way: with respect to any one of the indices A, fi, ... , and one of p, a, ... . In addition, the scalar (tensor of rank 0) does not appear, reducing to unity by virtue of the condition (59.6). §60. Time reversal and Kramers' theorem The symmetry of motion with respect to a change in the sign of the time is expressed in quantum mechanics by the fact that, if «/r is the wave function of a stationary state of the system, the "time-reversed" wave function (which we denote by «/i rev ) describes a possible state with the same energy. At the end of §18 it has been pointed out that */r rev is the same as the complex conjugate function </r*. In this simple form the statement applies to wave functions where the spin of particles is neglected. When spin is present, a refinement is necessary. Let us take the wave function of a particle of spin s in the form of the contra- variant spinor A / t — (of rank 2s). On taking the complex conjugate function ^,A/*...# we obtain a set of quantities which are transformed as components of a covariant spinor. Hence the operation of time reversal corresponds to a change from the wave function ifi^-- to a new wave function whose covariant com- ponents are given by .rev 0a,... = «A^- # . (60.1) For a given set of values of the indices A, /x, ... , the components of covariant and contravariant spinors correspond to values of the angular-momentum component which differ in sign. In terms of the functions «/r So ., therefore, time reversal corresponds to a change from ifj S(T to «/f S ,_ ff , as it should, since a change in the sign of the time changes the direction of the angular momentum. The exact relation is given by (60.1): t When these quantities are given, so are the mean values of the components of the vector s and all their powers and products 2, 3, ..., 2s at a time, which do not reduce to lower powers (see §55, Problem 3). §60 Time reversal and Kramers' theorem 207 In other words, the operation of time reversal requires the change ^->^,-.(-l) s -*- ( 60 - 2 ) Let us consider an arbitrary system of interacting particles. The orbital and spin angular momenta of such a system are not in general separately conserved when relativistic interactions are taken into account. Only the total angular momentum J is conserved. If there is no external field, each energy level of the system has (27+l)-fold degeneracy. When an external field is applied, the degeneracy is removed. The question arises whether the degeneracy can be removed completely, i.e. so that the system has only simple levels. This is closely related to the symmetry with respect to time reversal. In classical electrodynamics the equations are invariant with respect to a change in the sign of the time, if the electric field is left unchanged and the sign of the magnetic field is reversed.! This fundamental property of motion must be preserved in quantum mechanics. Hence, not only in a closed system but in any external electric field (there being no magnetic field), there is symmetry with respect to time reversal. The wave functions of the system are spinors xp^—, whose rank n is twice the sum of the spins s a of all the particles (n = 2 2 s a ); this sum may not be equal to the total spin S of the system. According to what was said above, we can assert that, in any electric field, the wave function and its time reversal must correspond to states with the same energy. If a level is non-degenerate, it is necessary that these states should be identical, i.e. the corresponding wave functions must be the same apart from a constant factor (both, of course, being expressed as similar (covariant or contra- variant) spinors). We write ^ v = C0 A/l ... or, by (60.1), ^...* = ClflX[i f (60.3) where C is a constant. Taking the complex conjugate of both sides of this equation, we obtain ^... = C *0 V *. We lower the indices on the left-hand side of the equation and correspond- ingly raise them on the right. This means that we multiply both sides of the equation by gaXgpn ••• an ^ sum over the indices A, /*, ... ; on the right-hand side we must use the fact that BcSJBfy. •••=(- l)V a ^-- As a result we have <Aam~. = C*(-i) n <A v "*. t See, for example, The Classical Theory of Fields, §17, and the end of §110 below. 208 Spin §60 Substituting 0V-* from (60.3), we find ^.. =(-i) n cc*^. This equation must be satisfied identically, i.e. we must have (— 1) W CC* = 1. Since, however, \C\ 2 is always positive, it is clear that this is possible only for even n (i.e. for integral values of the sum 2 s a ). For odd n (half-integral values of S s a ) the condition (60.3) cannot be fulfilled.f Thus we reach the result that an electric field can completely remove the degeneracy only for a system with an integral value of the sum of the spins of the particles. For a system with a half-integral value of this sum, in an arbitrary electric field, all the levels must be doubly degenerate, and complex conjugate spinors correspond to two different states with the same energy! (H. A. Kramers 1930). One further, mathematical, comment may be made. A relation of the form (60.3) with a real constant C is mathematically the condition that the components of the spinor may be put in correspondence with a set of real quantities, and may be called the condition for the spinor to be "real".|| The impossibility of fulfilling the condition (60.3) for odd n signifies that no real quantity can correspond to a spinor of odd rank. For even n, on the other hand, the condition (60.3) can be satisfied, and C can be real. In particular, a real vector can correspond to a symmetrical spinor of rank two if the condition (60.3) is satisfied with C = 1 : <A A "* = K (as is easily seen by means of (58.2) and (58.3)). The condition (60.3) with C = 1 is in fact the condition for a symmetrical spinor of any even rank to be "real". t When the sum S s„ is integral (or half-integral), all possible values of the total spin S of the system are also integral (or half-integral). X If the electric field possesses a high (cubic) symmetry, fourfold degeneracy may occur (see §99, including the Problem). || It is meaningless to call the spinor real in the literal sense, since complex conjugate spinors have different laws of transformation. CHAPTER IX IDENTITY OF PARTICLES §61. The principle of indistinguishability of similar particles In classical mechanics, identical particles (electrons, say) do not lose their "individuality", despite the identity of their physical properties. For we can imagine the particles at some instant to be "numbered", and follow the subsequent motion of each of these in its path ; then at any instant the particles can be identified. In quantum mechanics the situation is entirely different, as follows at once from the uncertainty principle. We have already mentioned several times that, by virtue of the uncertainty principle, the concept of the path of an electron ceases to have any meaning. If the position of an electron is exactly known at a given instant, its co-ordinates have no definite values even at an infinitely close subsequent instant. Hence, by localising and numbering the electrons at some instant, we make no progress towards identifying them at subsequent instants ; if we localise one of the electrons, at some other instant, at some point in space, we cannot say which of the electrons has arrived at this point. Thus, in quantum mechanics, there is in principle no possibility of separ- ately following each of a number of similar particles and thereby distinguish- ing them. We may say that, in quantum mechanics, identical particles entirely lose their "individuality". The identity of the particles with respect to their physical properties is here very far-reaching: it results in the complete indistinguishability of the particles. This principle of the indistinguishability of similar particles, as it is called, plays a fundamental part in the quantum-mechanical investigation of systems composed of identical particles. Let us start by considering a system of only two particles. Because of the identity of the particles, the states of the system obtained from each other by merely interchanging the two particles must be completely equivalent physically. This means that, as a result of this inter- change, the wave function of the system can change only by an unimportant phase factor. Let ip(g lf | 2 ) be the wave function of the system, £ x and £ 2 con- ventionally denoting the three co-ordinates and the spin projection for each particle. Then we must have where a is some real constant. By repeating the interchange, we return to the original state, while the function is multiplied by e 2toc . Hence it follows that e %i(X = 1, or e ia = ±1. Thus *&.&)= ±*(&,*i). 209 210 Identity of Particles §61 We thus reach the result that there are only two possibilities: the wave function is either symmetrical (i.e. it is unchanged when the particles are inter- changed) or antisymmetrical (i.e. it changes sign when this interchange is made). It is obvious that the wave functions of all the states of a given system must have the same symmetry; otherwise, the wave function of a state which was a superposition of states of different symmetry would be neither sym- metrical nor antisymmetrical. This result can be immediately generalised to systems consisting of any number of identical particles. For it is clear from the identity of the particles that, if any pair of them has the property of being described by, say, sym- metrical wave functions, any other pair of such particles has the same pro- perty. Hence the wave function of identical particles must either be un- changed when any pair of particles are interchanged (and hence when the particles are permuted in any manner), or change sign when any pair are interchanged. In the first case we speak of a symmetrical wave function, and in the second case of an antisymmetrical one. The property of being described by symmetrical or antisymmetrical wave functions depends on the nature of the particles. Particles described by antisymmetrical functions are said to obey Fermi-Dirac statistics (or to be fermions), while those which are described by symmetrical functions are said to obey Bose-Einstein statistics (or to be bosons).^ Relativistic quantum mechanics shows that the statistics obeyed by particles is uniquely related to their spin : particles with half- integral spin are fermions, and those with integral spin are bosons. The statistics of complex particles is determined by the parity of the number of elementary fermions entering into their composition. For an interchange of two identical complex particles is equivalent to the simul- taneous interchange of several pairs of identical elementary particles. The interchange of bosons does not change the wave function, while the inter- change of fermions changes its sign. Hence complex particles containing an odd number of elementary fermions obey Fermi statistics, while those containing an even number obey Bose statistics. This result is, of course, in agreement with the above rule, since a complex particle has an integral or a half-integral spin according as the number of particles with half-integral spin entering into its composition is even or odd. Thus atomic nuclei of odd atomic weight (i.e. containing an odd number of neutrons and protons) obey Fermi statistics, and those of even atomic weight obey Bose statistics. For atoms, which contain both nuclei and electrons, the statistics is evidently determined by the parity of the difference between the atomic weight and the atomic number. t This terminology refers to the statistics which describes a perfect gas composed of particles with antisymmetrical and symmetrical wave functions respectively. In actual fact we are concerned here not only with a different statistics, but essentially with a different mechanics. Fermi statistics was proposed by E. Fermi for electrons in 1926, and its relation to quantum mechanics was elucidated by P. A. M. Dirac (1926). Bose statistics was proposed by S. N. Bose for light quanta, and generalised by A. Einstein (1924). §61 The principle of indistinguishability of similar particles 211 Let us consider a system composed of N identical particles, whose mutual interaction can be neglected. Let «/» 1} i/r 2 , ... be the wave functions of the vari- ous stationary states which each of the particles separately may occupy, f The state of the system as a whole can be denned by giving the numbers of the states which the individual particles occupy. The question arises how the wave function Y of the whole system should be constructed from the functions if/ ly ift 2 , ... . Let Px,p 2 , ... ,p N be the numbers of the states occupied by the individual particles (some of these numbers may be the same). For a system of bosons, the wave function T(£i, £ 2 , ..., $n) is given by a sum of products of the form 1^(6)*, (*•>•" **<**>' with all possible permutations of the different suffixes p lt p 2 , ... ; this sum clearly possesses the required symmetry property. Thus, for example, for a system of two particles T&.&) = [&> (*iW, (&)+*„ ($M>p &)]/V2; (61.1) we suppose that^ # p 2 . The factor \\y/2 is introduced for normalisation purposes; all the functions iff lt 2 , ... are orthogonal and are supposed normal- ised. For a system of fermions, the wave function *F is an antisymmetrical combination of these products. It can be written in the form of a deter- minant iM6) M&) Y = VNl l^tfi) M« iM&) *..(&) ^ (for) i <Ap 2 (^) (61.2) Here an interchange of two particles corresponds to an interchange of two columns of the determinant, as a result of which the latter, as is well known, changes sign. For a system composed of two particles we have y = [M^M^-MtoMfiWa. (61.3) The following important result is a consequence of the expression (61.2). If among the numbers p lf p 2 ,... any two are the same, two rows of the determinant are the same, and it therefore vanishes identically. It will be different from zero only when all the numbers p lt p 2 , ... are different. Thus, in a system consisting of identical fermions, no two (or more) particles can be in the same state at the same time. This is called PaulV s principle (1925). t If there is a strong interaction between the particles we cannot, of course, speak of such states. 212 Identity of Particles §62 §62. Exchange interaction The fact that Schrodinger's equation does not take account of the spin of particles does not invalidate this equation or the results obtained by means of it. This is because the electrical interaction of the particles does not depend on their spins.f Mathematically, this means that the Hamiltonian of a system of electrically interacting particles (in the absence of a magnetic field) does not contain the spin operators, and hence, when it is applied to the wave function, it has no effect on the spin variables. Hence Schrodinger's equation is actually satisfied by each component of the wave function; in other words, the wave function ^(r^ a x \ r 2 , <r 2 ; ...) of the system of particles can be written in the form of a product x(°i> ff a»— W( r i> r »— ) of a function <j> of the co-ordinates of the particles only and a function x of the spins. We call the former a co-ordinate or orbital wave function, and the latter a spin wave function. Schrodinger's equation essentially determines only the co-ordinate function <f>, the function x remaining arbitrary. In any instance where we are not interested in the actual spin of the particles, we can therefore use Schrodinger's equation and regard as the wave function the co-ordinate function alone, as we have done hitherto. However, despite the fact that the electrical interaction of the particles is independent of their spin, there is a peculiar dependence of the energy of the system on its total spin, arising ultimately from the principle of indistinguishability of similar particles. Let us consider a system consisting of only two identical particles. By solving Schrodinger's equation we find a series of energy levels, to each of which there corresponds a definite symmetrical or antisymmetrical co- ordinate wave function </>(ri, r 2 ). For, by virtue of the identity of the particles, the Hamiltonian (and therefore the Schrodinger's equation) of the system is invariant with respect to interchange of the particles. If the energy levels are not degenerate, the function <f>(ri, r 2 ) can change only by a constant factor when the co-ordinates ri and r 2 are interchanged; repeating this interchange, we see that this factor can only be J ± 1. Let us first suppose that the particles have zero spin. The spin factor for such particles is absent altogether, and the wave function reduces to the co-ordinate function ^(r 1? r 2 ), which must be symmetrical (since particles with zero spin obey Bose statistics). Thus not all the energy levels obtained by a formal solution of Schrodinger's equation can actually exist; those to which antisymmetrical functions <j> correspond are not possible for the system under consideration. t This is true only so long as we consider the non-relativistic approximation. When relativistic effects are taken into account, the interaction of charged particles does depend on their spin. J When there is degeneracy we can always choose linear combinations of the functions belonging to a given level, such that this condition is again satisfied. §62 Exchange interaction 213 The interchange of two similar particles is equivalent to the operation of inversion of the co-ordinate system (the origin being taken to bisect the line joining the two particles). On the other hand, the result of inversion is to multiply the wave function <j> by ( — 1)*, where / is the orbital angular momen- tum of the relative motion of the two particles (see §30). By comparing these considerations with those given above, we conclude that a system of two identical particles with zero spin can have only an even orbital angular mo- mentum. Next, let us suppose that the system consists of two particles with spin £ (say, electrons). Then the complete wave function of the system (i.e. the product of the function ^(r 1} r 2 ) and the spin function x(<*v cr 2 )) must certainly be antisymmetrical with respect to an interchange of the two electrons. Hence, if the co-ordinate function is symmetrical, the spin function must be antisymmetrical, and -vice versa. We shall write the spin function in spinor form, i.e. as a spinor x V of rank two, each of whose indices corresponds to the spin of one of the electrons. A symmetrical spinor (^V = x^) corre- sponds to a function symmetrical with respect to the spins of the two particles, and an antisymmetrical spinor (x^ = — x^) to an antisymmetrical func- tion. We know, however, that a symmetrical spinor of rank two describes a system with total spin unity, while an antisymmetrical spinor reduces to a scalar, corresponding to zero spin. Thus we reach the following conclusion. The energy levels to which there correspond symmetrical solutions ^(r 1? r 2 ) of Schrodinger's equation can actually occur when the total spin of the system is zero, i.e. when the spins of the two electrons are "antiparallel", giving a sum of zero. The values of the energy belonging to antisymmetrical functions (frfa, r 2 ), on the other hand, require a value of unity for the total spin, i.e. the spins of the two electrons must be "parallel". In other words, the possible values of the energy of a system of electrons depend on their total spin. For this reason we can speak of a peculiar inter- action of the particles which results in this dependence. This is called exchange interaction. It is a purely quantum effect, which entirely vanishes (like the spin itself) in the passage to the limit of classical mechanics. The following situation is characteristic of the case of a system of two electrons which we have discussed. To each energy level there corresponds one definite value of the total spin, or 1. This one-to-one correspondence be- tween the spin values and the energy levels is preserved, as we shall see below (§63), in systems containing any number of electrons. It does not hold, however, for systems composed of particles whose spin exceeds \. Let us consider a system of two particles, each with arbitrary spin s. Its spin wave function is a spinor of rank 4s: Aji... po... *~28 IT"' half (2s) of whose indices correspond to the spin of one particle, and the other 214 Identity of Particles §62 half to that of the other particle. The spinor is symmetrical with respect to the indices in each group. An interchange of the two particles corresponds to an interchange of all the indices A, [i, ... of the first group with the indices p, a, ... of the second group. In order to obtain the spin function of a state of the system with total spin S, we must contract this spinor with respect to 2s— S pairs of indices (each pair containing one index from X, /u, ... and one from p, a, ...), and symmetrise it with respect to the remainder; as a result we obtain a symmetrical spinor of rank 2S. However, the contraction of a spinor with respect to a pair of indices means, as we know, the construction of a combination antisymmetrical with respect to these indices. Hence, when the particles are interchanged, the spin wave function is multiplied by(-l)»-« On the other hand, the complete wave function of a system of two particles must be multiplied by ( — l) 2s when they are interchanged (i.e. by +1 for integral s and by — 1 for half-integral s). Hence it follows that the symmetry of the co-ordinate wave function with respect to an interchange of the particles is given by the factor ( — l) s , which depends only on S. Thus we reach the result that the co-ordinate wave function of a system of two identical particles is symmetrical when the total spin is even, and antisymmetrical when it is odd. Recalling what was said above concerning the relation between interchange of the particles and inversion of the co-ordinate system, we conclude also that, when the spin S is even (odd), the system can have only an even (odd) orbital angular momentum. We see that here also a certain dependence is revealed between the possible values of the energy of the system and the total spin, but this dependence is not necessarily one-to-one. The energy levels to which there correspond symmetrical (antisymmetrical) co-ordinate wave functions can occur for any even (odd) value of S. Let us calculate how many different states of the system there are with even and odd S. The quantity S takes 2s +1 values: 2s, 2s— 1, ..., 0. For any given S there are 2S+1 states differing in the value of the s-component of the spin ((2s+ 1) 2 different states altogether). Let s be integral. Then, among the 2s +1 values of S, s+1 are even and s odd. The total number of states with even S is equal to the sum s iUF +1 > =(2s+1)(s+1): the remaining s(2s+l) states have odd S. Similarly, we find that, when s is half-integral, there are s(2s+l) states with even values of S and (s+l)(2s+l) with odd values. PROBLEMS Problem 1. Determine the exchange splitting of the energy levels of a system of two electrons, regarding the interaction of the electrons as a perturbation. Solution. Let the particles be (when their interaction is neglected) in states with orbital §62 Exchange interaction 215 wave functions <i>i(r) and fa(r). The states of the system with total spin S = and S = 1 correspond to symmetrised and antisymmetrised products respectively: 1 <f> = —dM r i)M r 2)±M r 2)M r i)l The mean value of the operator of the interaction U(rz— ri) of the particles in these states is A ± J, where A = J j UlMriTlU^WdVidVz, J= \ \ C/0i(ri)^i*(r 2 )^2(r2)^2*(ri) dFi dV 2 , the latter being called the exchange integral. Omitting the additive constant A, which is not an exchange term, we therefore find the level shifts Aj£ = J, Aj?i = —J (where the suffix indicates the value of S). These quantities can be represented as the eigenvalues of the spin "exchange operator" t Pexch= -17(1+48!. s 2 ); (1) the eigenvalues of the product si . S2 are derived in §55, Problem 2. If the electrons belong to different atoms, for example, the exchange integral decreases exponentially with increasing distance R between the atoms. It is clear from the form of the integrand that this integral is determined by the "overlap" of the wave functions of the states <f>(ri) and ^2(1*2) ; using the asymptotic law of decrease of the wave functions of states of a discrete spectrum (cf. (21.6)), we find that J „ r-lKi+icm, K1 = V(2m|£i|)/£, K2 = ^/(2m\E 2 \)jh, where Ei and E2 are the energy levels of the electron in the two atoms. Problem 2. The same as Problem 1, but for a system of three electrons. Solution. Using formula (1), Problem 1, we can write the operator of pairwise exchange interaction in a system of three electrons as ^exch = — 2 J a& (^ + 2s a .S 6 ), (1) where the summation is over pairs of particles 12, 13 and 23. The matrix elements of the operators s a . s& between states with different values of the pair of numbers o a> at, are given by formulae (55.6) as 11 1 — i i 1 (Sa.S&) u = 4, (Sa.Sftji,-! = — 4, ( s a« s fc)-i,i = 2 . We first determine the energy corresponding to the greatest possible value of the total-spin component Ms = 01+02 + 03, viz. Ms = 3/2. This gives the energy of the state with total spin »S = 3/2. On calculating the corresponding diagonal matrix element of the operator (1), we find A-E3/2 = — (7i2 + /l3 + /23)- Next we take states with Ms = i. This value can occur in three ways, depending on which of the numbers 01, 02, 031s —$ (the other two being $). Thus for these states we should have a secular equation of the third degree. The calculation can, however, be simplified immediately by noting that one of the roots of this equation must correspond to the energy already found for the state with 5 = 3/2, and the secular equation must therefore have the factor AE—AE3/2. In this way the calculation of the free term in the cubic equation can be avoided. J t First used by Dirac. } This device is particularly useful in similar calculations for systems with a larger number of particles . 216 Identity of Particles §63 The leading terms of the equation are found to be (A£)3 + (7i2 + /l3 + /23)(A£ 2 )+[/i27l3 + A2/23 + A3^3- -9(7i2 2 + 7l3 2 + ^3 2 )/4]AE+ ... = 0. Dividing by A£ '+ J12 + J13+ J23, we find the two energy levels corresponding to states with spins S = i: AEl/ 2 = ±[(/l2 2 + /l3 2 + ^23 2 )-/l2^13-il2i23-Jl3i23] 1/2 . Thus there are three energy levels, in accordance with the calculation in §63, Problem. Problem 3. In which states can the Be 8 nucleus decay into two oe-particles? Solution. Since the a-particle has no spin, a system of two a-particles can only have an even orbital angular momentum (equal to the total angular momentum), and its states are even. The decay in question is therefore possible only from even states of the Be 8 nucleus with even total angular momentum. §63. Symmetry with respect to interchange By considering a system composed of only two particles, we have been able to show that its co-ordinate wave functions <£(r x , r 2 ) for the stationary states must be either symmetrical or antisymmetrical. In the general case of a sys- tem of an arbitrary number of particles, the solutions of Schrodinger's equation (the co-ordinate wave functions) need not necessarily be either sym- metrical or antisymmetrical with respect to the interchange of any pair of particles, as the complete wave functions (which include the spin factor) must be. This is because an interchange of only the co-ordinates of two par- ticles does not correspond to a physical interchange of them. The physical identity of the particles here leads only to the fact that the Hamiltonian of the system is invariant with respect to the interchange of the particles, and hence, if some function is a solution of Schrodinger's equation, the functions ob- tained from it by various interchanges of the variables will also be solutions. Let us first of all make some remarks regarding interchanges in general. In a system of N particles, N\ different permutations in all are possible. If we imagine all the particles to be numbered, each permutation can be represented by a definite sequence of the numbers 1, 2, 3, ... . Every such sequence can be obtained from the natural sequence 1, 2, 3, ... by successive interchanges of pairs of particles. The permutation is called even or odd, according as it is brought about by an even or odd number of such inter- changes. We denote by P the operators of permutations of N particles, and introduce a quantity S P which is +1 if P is an even permutation and —1 if it is odd. If <f> is a function symmetrical with respect to all the particles, we have while, if <f> is antisymmetrical with respect to all the particles, then P<f> = bp4>. §63 Symmetry zvith respect to interchange 217 From an arbitrary function ^(r^ r 2 , ... , r^), we can form a symmetrical function by the operation of symmetrisation, which can be written <f> sym = constant xS^, (63.1) where the summation extends over all possible permutations. The formation of an antisymmetrical function (an operation sometimes called alternation) can be written as 0^ = constant x S 8 P P<f>. (63.2) Let us return to considering the behaviour, with respect to permutations, of the wave functions ^ of a system of identical particles, j- The fact that the Hamiltonian i? of the system is symmetrical with respect to all the particles means, mathematically, that jff commutes with all the permutation operators P. These operators, however, do not commute with one another, and so they cannot be simultaneously brought into diagonal form. This means that the wave functions^ cannot be so chosen that each of them is either symmetri- cal or antisymmetrical with respect to all interchanges separately.]; Let us try to determine the possible types of symmetry of the functions <f>{r x , r 2 , ... , r N ) of N variables (or of sets of several such functions) with respect to permutations of the variables. The symmetry must be such that it "cannot be increased", i.e. such that any additional operation of symmetri- sation or alternation, on being applied to these functions, would reduce them either to linear combinations of themselves or to zero identically. We already know two operations which give functions with the greatest possible symmetry: symmetrisation with respect to all the variables, and alternation with respect to all the variables. These operations can be general- ised as follows. We divide the set of all the N variables r 1} r 2 , ... , r^ (or, what is the same thing, the suffixes 1, 2, 3, ... , N) into several sets, containing N ly N 2 , ... ele- ments (variables); N 1 +N 2 + ... = N. This division can be conveniently shown by a diagram (known as a Young diagram) in which each of the num- bers N lt N 2 , ... is represented by a line of several cells (thus, Fig. 21 gives a diagram of the divisions 6+4+4+3+3 + 1+1 and 7+5+5+3 + 1 + 1 for N = 22); one of the numbers 1, 2, 3, ... is to be placed in each square. If we place the lines in order of decreasing length (as in Fig. 21), the diagram contains not only successive horizontal rows, but also vertical columns. Let us symmetrise an arbitrary function ^(r x , r 2 , ... , r N ) with respect to the variables in each row. The alternation operation can then be performed only with respect to the variables in different rows; alternation with respect to a pair of variables in the same row clearly gives zero identically. t From the mathematical point of view, the problem is to find irreducible representations of the permutation group. A detailed account of the mathematical theory of permutation (or symmetry) groups is given by H. Weyl, The Theory of Groups and Quantum Mechanics, Methuen, London 1931 ; D. E. Rutherford, Substitutional Analysis, University Press, Edinburgh 1948; F. D. Murnaghan, The Theory of Group Representations, Johns Hopkins Press, Baltimore 1938. J Except for a system of only two particles, where there is a single interchange operator, which can be brought into diagonal form simultaneously with ff. 218 Identity of Particles §63 Fig. 21 Having chosen one variable from each row, we can, without loss of gener- ality, regard them as being in the first cells in each row (after symmetrisation, the order of the variables among the cells in each row is immaterial) ; let us alternate with respect to these variables. Having then deleted the first column, we alternate with respect to variables chosen one from each row in the thus "curtailed" diagram; these variables can again be regarded as being in the first cells of the "curtailed" rows. Continuing this process, we finally have the function first symmetrised with respect to the variables in each row and then alternated with respect to the variables in each column. After alternation, of course, the function in general ceases to be symmetrical with respect to the variables in each row. The symmetry is preserved only with respect to the variables in the cells of the first row which project beyond the other rows. Having distributed the N variables in various ways among the rows of a Young diagram (the distribution among the cells in each row is immaterial), we thus obtain a series of functions, which are transformed linearly into one another when the variables are permuted in any manner, f However, it must be emphasised that not all these functions are linearly independent ; the number of independent functions is in general less than the number of possible distri- butions of the variables among the rows of the diagram. We shall not pause here, however, to discuss this more closely. % Thus any Young diagram determines some type of symmetry of functions with respect to permutations. By constructing all the possible Young dia- grams (for a given N), we find all possible types of symmetry. This amounts to dividing the number N in all possible ways into a sum of smaller terms, including the number N itself; thus for N = 4 the possible partitions are 4,3 + 1,2+2,2+1 + 1,1 + 1 + 1 + 1. To each energy level of the system we can make correspond a Young dia- gram which determines the permutational symmetry of the appropriate solutions of Schrodinger's equation; in general, several different functions correspond to each value of the energy, and these are transformed linearly into f It would be possible to perform the symmetrisation and alternation in the reverse order: to alter- nate with respect to the variables in each column, and then to symmetrise with respect to those in the rows. This, however, would give effectively the same thing, since the functions obtained by the two methods are linear combinations of one another. J For particles with spin i the number of independent functions (i.e. the dimension of the irreducible representation) is derived in the Problem below. §63 Symmetry with respect to interchange 219 each other by permutations.! However, it must be emphasised that this does not signify any additional physical degeneracy of the energy levels. All these different co-ordinate wave functions, multiplied by the spin functions, enter into a single definite combination — the complete wave function — which satisfies (according to the number of particles) the condition of symmetry or antisymmetry. Among the various types of symmetry there are always (for any given N) two to each of which only one function corresponds. One of these cor- responds to a function symmetrical with respect to all the variables, and the other to one which is similarly antisymmetrical ; in the first case, the Young diagram consists of a single row of N cells, and in the second case of a single column. Let us now consider the spin wave functions x(cr l5 tr 2 , ... , a N ). Their kinds of symmetry with respect to permutations of the particles are given by the same Young diagrams, with the components of the spins of the particles taking the part of variables. There arises the question of what diagram must correspond to the spin function for a given diagram of the co-ordinate func- tion. Let us first suppose that the spin of the particles is integral. Then the complete wave function if/ must be symmetrical with respect to all the particles. For this to be so, the symmetry of the spin and co-ordinate functions must be given by the same Young diagram, and the complete wave function ijj is expressed as definite bilinear combinations of the two ; we shall not here pause to examine more closely the problem of constructing these combinations. Next, suppose the spin of the particles to be half-integral. Then the com- plete wave function must be antisymmetrical with respect to all the particles. It can be shown that, for this to be so, the Young diagrams for the co-ordinate and spin functions must be obtained from each other by interchanging rows and columns (as in the two diagrams shown in Fig. 21). Let us consider in more detaii the important case of particles with spin \ (electrons, for instance). Each of the spin variables ct 1? o- 2 , ... here takes only the two values ±|. Since a function antisymmetrical with respect to any two variables vanishes when these variables take the same value, it is clear that the function x can be alternated only with respect to pairs of variables ; if we alternate with respect to even three variables, two of them must always take the same value, so that we have zero identically. Thus, for a system of electrons, the Young diagrams for the spin functions can contain columns of only one or two cells (i.e. only one or two rows) ; in the Young diagrams for the co-ordinate functions, the same is true of the number of columns. The number of possible types of permutational sym- metry for a system of N electrons is therefore equal to the number of possible partitions of the number N into a sum of ones and twos. When N is even, this number is %N+1 (partitions with 0, 1, ... , %N twos), while if N is odd t The existence of this "permutational degeneracy" is related to the fact that the permutation oper- ators commuting with the Hamiltonian do not in general commute with one another (see the middle of §10). 220 Identity of Particles §63 it is f(iV+l) (partitions with 0, 1, ... , \{N— 1) twos). Thus, for instance, Fig. 22 shows the possible Young diagrams (co-ordinate and spin) for N = 4. It is easy to see that each of these types of symmetry (i.e. each of the Young diagrams) corresponds to a definite total spin S of the system of electrons. We shall consider the spin functions in spinor form, i.e. as spinors ^A/x... f ran k JV, whose indices (each of which corresponds to the spin of an individual particle) will be the variables that are arranged in the cells of the Young diagrams. Let us examine the Young diagram consisting of two rows with N x and N % cells {N x +N% = N, and N x > iV 2 ). In each of the first N 2 columns there are two cells, and the spinor must be antisymmetrical with respect to the corresponding pairs of indices. With respect to the indices in the last n = N 1 —N. 3 ells in the first row, however, it must be symmetrical. As we know, such a spinor of rank N reduces to a symmetrical spinor of rank n, to which there corresponds a total spin S = \n. Returning to the Young diagrams for the co-ordinate functions, we can say that the diagram with n rows each of one cell corresponds to a total spin S = \n. For even N, the total spin can take integral values from to |iV, while for odd N it can take half-integral values from | to $N, as it should. We emphasise that this one-to-one correspondence between the Young diagrams and the total spin holds only for systems of particles with spin | ; we have seen this, for a system of two particles, in the previous section. f PROBLEM Determine the number of energy levels with different values of the total spin S, for a system of N particles of spin J. Solution. A given value of the projection of the total spin of the system, M s = 2 a, can be obtained in f(M s ) = N\l(W+M s )\(iN-M s )\ ways; this is the number of combinations of N elements £N+M S at a time, since we put a = £ for %N+M S particles and a = —\ for the remainder. To each energy level with a given S, there correspond 2S+1 states with values M s = S, S—l, ... , —S of the projected spin. Hence it is easy to see that the number of different energy levels with a given value of S is n(S) =f(S)-f(S+l) =Nl (2S+ l)l(iN+ S+1)\QN-S)\ f In the third footnote to §20, we have remarked that, for a system of several identical particles, we cannot assert that the wave function of the stationary state of lowest energy is without nodes. We shall now amplify this statement and elucidate its origin. The wave function (that is, the co-ordinate function), if it has no nodes, must certainly be symmetri- cal with respect to all the particles; for, if it were antisymmetrical with respect to the interchange of any pair of particles 1, 2, it would vanish for r x = r 2 . If, however, the system consists of three or more electrons, no completely symmetrical co-ordinate wave function is possible : the Young diagram of the co-ordinate function cannot have rows with more than two cells. Thus, although the solution of Schrodinger's equation which corresponds to the lowest eigenvalue is without nodes (by the theorem of the variational calculus), this solution may be physically inadmis- sible; the smallest eigenvalue of Schrodinger's equation will not then correspond to the normal state of the system, and the wave function of this state will in general have nodes. For particles with a half-integral spin s, this situation occurs in systems with more than 2s +1 particles. For systems of bosons, a completely symmetrical co-ordinate wave function is always possible. §64 Second quantisation. The case of Bose statistics 221 J S=2 S = 1 Fig. 22 SrQ The total number of different energy levels is n = Sn(5)=/(0)=AT!/[(|iV)!p s for even N, and »=/(i)=M/(*tf +*)!(*#-*)' for odd JV. §64. Second quantisation. The case of Bose statistics In the quantum-mechanical investigation of systems consisting of a very large number of identical particles, interacting in any manner, there is a useful method of considering the problem, known as second quantisation. This method is necessary in relativistic theory also, where we have to deal with systems in which the number of particles is itself variable^ Let us first consider systems of bosons. We denote by 0i(£), ^2(£)> ••• some complete set of orthogonal and normal- ised wave functions. These may, for instance, correspond to the stationary states of a single particle in some external field. We emphasise that the choice of this field is arbitrary ; it need not be the same as the actual field acting on the particles in the physical system considered. As in §61, | denotes the assembly of the co-ordinates and spin projection a of a particle. Let us consider, in a purely formal manner, a system of N non-interacting particles, in the field selected. Then every particle is in one of the states «/f 1} ?/f 2 , ... . Let N t be the number of particles in the state ^; it may, of course, be zero. If the numbers N±, N%, ... are given (clearly 2 Nt = N), the state of the system as a whole is determined; we shall indicate these numbers by suffixes to the wave function M^n,... °f tne system. Let us seek to construct a mathematical formalism in which the occupation numbers N lt N 2 , ... of the states (and not the co-ordinates of the particles) play the part of independent variables. f The method of second quantisation was developed by P. A. M. Dirac (1927) for particles obeying Bose statistics, and later extended to Fermi particles by E. Wigner and P. Jordan (1928). 222 Identity of Particles §64 The function ^¥ NiNz _ is a symmetrised (the particles obeying Bose statistics) sum of products of the functions fa. Let us write it in the form Yur v... = VWNJ ...INI) S fa (Qfa (&) ... fa (fr). (64.1) i. «s 1 2 N - Here p ly p 2 , ... , ^v are the ordinal numbers of the states in which the indivi- dual particles are, and the sum is taken over all permutations of those suffixes Pi>p2> ••• >Pn which are different. The numbers A^ show how many of the suffixes p x , p 2 , ... , p N have the value i. The total number of terms in the sum (64.1) is evidently Nl/NjlNJi.... The constant factor in (64.1) is chosen so that the function is normalised; by virtue of the orthogonality of the functions fa, on integrating! tne square IT/v^... ! 2 with respect to | x , £ 2 , ... , g N all the terms vanish except the squared modulus of each term in the sum. Next, let / (1) a be the operator of some physical quantity pertaining to the <zth particle, i.e. acting only on functions of £ a . We introduce the operator ^ (1) = S/(« a , (64.2) which is symmetrical with respect to all the particles (the summation being over all particles), and determine its matrix elements with respect to the wave functions (64.1). First of all, it is easy to see that the matrix elements will be different from zero only for transitions which leave the numbers N ± , N 2 , ... unchanged (diagonal elements) and for transitions where one of these numbers is increased, and another decreased, by unity. For, since each of the operators /<!>„ acts only on one function in the product fajiijfa^) ... fa^), its matrix elements can be different from zero only for transitions whereby the state of a single particle is changed ; this, however, means that the number of particles in one state is diminished by unity, while the number in another state is correspondingly increased. The calculation of these matrix elements is in principle very simple ; it is easier to do it oneself than to follow an account of it. Hence we shall give only the result of this calculation. The non- diagonal elements are Fll) N^N k =f a) ikV(NiN k ). (64.3) We shall indicate only those suffixes with respect to which the matrix element is non-diagonal, omitting the remainder for brevity. Here/ (1) <fc is the matrix element f w nc=jfa*(t)f w fa(t)dt. (64.4) f By integration over £ we conventionally understand integration over the co-ordinates and sum- mation over a. §64 Second quantisation. The case of Bose statistics 223 It must be borne in mind that the operators / (1) differ only in the naming of the variables on which they act, and hence the integrals / (1) i& are indepen- dent of a. The diagonal matrix elements of F (1) are the mean values of the quantity FV> in the states Yj^jy,.... Calculation gives F» = ?/««#<• (64.5) i We now introduce the operators d iy which play a leading part in the method of second quantisation; they act, not on functions of the co-ordinates, but on the variables N lt N it ... , and are defined as follows. When acting on the function Y# # ., the operator d t decreases the suffix N t by unity, and at the same time it multiplies the wave function by ^N t : d^ Nl N2 ... Ni ... = V% n 2 .... *hu. (64.6) We can say that the operator a\ diminishes by one the number of particles in the ith state; it is called an annihilation operator. It can be represented in the form of a matrix whose only non-zero element is (a,VV* = VNi. (64.7) The operator &f which is the Hermitian conjugate of d i is, by definition (see §3), represented by a matrix with an element i.e. ta + #-i =VNi ' (64 ' 8) This means that, when acting on the function x i r NlNt ..., it increases the suffix N t by unity: d^¥ Nl n 2 ...N t ... = VW+Wn, n 2 ...,N i+1 (64.9) In other words, the operator di + increases by one the number of particles in the tth state, and is therefore called a creation operator. The product of operators d^d*, acting on the wave function, evidently multiplies it by a constant simply, leaving unchanged all the variables N lt N 2 , ... : the operator d t diminishes N t by unity, and d t + then restores it to its original value. Direct multiplication of the matrices (64.7) and (64.8) shows that afdi is represented, as we should expect, by a diagonal matrix whose diagonal elements are N t . We can write dt+dt = N t . (64.10) Similarly, we find that <W=iVVr-l. (64.11) 224 Identity of Particles §64 Hence the commutation rule for the operators d t and d t + is d i d i +—d i + d i = 1. (64.12) The operators d t and d k (or d t and d k + ) with i and k different act on different variables (N t and N k ), and of course commute: didic—dicdi = 0, didk + -dk + di = (i ^ A). (64.13) From the above properties of the operators d t , df it is easy to see that the operator /» = S/ (1) a^ + 4 (64.14) is the same as the operator (64.2). For all the matrix elements calculated from (64.7), (64.8) are the same as the elements (64.3), (64.5). This is a very important result. In formula (64. 14), the quantities/ Q-\ k are simply numbers. Thus we have been able to express an ordinary operator (of the form (64.2)), acting on functions of the co-ordinates, in the form of an operator acting on functions of new variables, the occupation numbersf N t . The result which we have obtained is easily generalised to operators of other forms. Let jft2) = S/(2) a6j (6415) a>o where / (2) a6 is the operator of a physical quantity pertaining to two particles at once, and hence acts on functions of g a and g b . Similar calculations show that this operator can be expressed in terms of the operators d t , d t + by where (f ( X = If ^(^*tfi)M(6)ao d&d*,. The matrices calculated for (64.15) and (64.16) are the same. The generalisa- tion of these formulae to operators of any other form symmetrical with respect to all the particles (of the form / (3) = S/ 3 > a6c etc.) is obvious. Finally, it remains to express, in terms of the operators d iy the Hamiltonian & of the physical system of N identical interacting particles that is actually being considered. The operator i? is, of course, symmetrical with respect to all the particles. In the non-relativistic approximation, J it is independent f Formula (64.14) bears a similarity to the expression (11.1) f = ^>fika*a k for the mean value of a quantity/, expressed in terms of the coefficients a { in the expansion of the wave function of a given state in terms of the wave functions of the stationary states. This is the reason for calling this method the second quantisation method. % In the absence of a magnetic field. §64 Second quantisation. The case of Bose statistics 225 of the spins of the particles, and can be represented in a general form as follows : i? = S^)+S^(W)+ s £/< 3 >(r a ,r 6 ,r c )+.... (64.17) a a>b a>b>c Here # (1) is the part of the Hamiltonian which depends on the co-ordinates of the ath particle only : #(D a = -(#72m)Aa+tf (1) (»-«), ( 64 - 18 ) where C/ (1) (r a ) is the potential energy of a single particle in the external field. The remaining terms in (64.17) correspond to the mutual interaction energy of the particles; for convenience, the terms depending on the co-ordinates of two, three, etc. particles have been separated. This representation of the Hamiltonian enables us to apply formulae (64.14), (64.16) and their analogues directly. Thus # = JflW^A+^+1 J^Xfl i +a?d 1 A+ ... • (64.19) This gives the required expression for the Hamiltonian in the form of an operator acting on functions of the occupation numbers. For a system of non-interacting particles, only the first term in the expres- sion (64.19) remains: i,k If the functions ^ are taken to be the eigenfunctions of the Hamiltonian #(*> of an individual particle, the matrix H (X \ k is diagonal, and its diagonal elements are the eigenvalues e i of the energy of the particle. Thus ft = S efdi+di', i replacing the operator a^ by its eigenvalues (64.10), we have for the energy levels of the system the expression E = 2 €i N it a trivial result which could have been foreseen. The formalism which we have developed can be put in a somewhat more compact form by introducing the operatorsf t Attention is drawn to the analogy between these expressions and the expansion T = S aWt of the wave function in terms of the eigenfunctions of an operator (cf . the third footnote to this sec- tion). 226 Identity of Particles §64 where the variables £ are regarded as parameters. By what has been said above concerning the operators d it df, it is clear that the operator *F de- creases the total number of particles in the system by one, while ¥+ increases it by one. It is easy to see that the operator ¥ + (f ) creates a particle at the point | . For the result of the action of the operator of is to create a particle in a state with wave function «/^(£). Hence it follows that the result of the action of the operator ^(lo) is to create a particle in a state with wave function S <Ai*(£)^(£o)> or ( bv the general formula (5.12)) with wave functionf S(£— | ), which corresponds to a particle with definite values of the co-ordinates (and spin). The commutation rules for ¥ and 4^+ are obtained at once from those for d it d t + . It is evident that *(*)¥(*' )-¥(*')*(0 = 0; (64.21) we also have ^)T+(f)-^+(n^) = 2&GrW(f), or n^ + (n-^ + (n^) = w-n (64.22) The expression (64.14) for the operator (64.2) can be written, using these new operators, in the form fr>=f¥+{f)fw*(QiX, (64.23) where it is understood that the operator / (1) acts on functions of the parameters | in ^"(l)- For, substituting (64.20), we have /» =^J4>i*(i)f w Ui)^.d^d k = s/ (1 W< + 4, which is the same as (64.14). Similarly, we have instead of (64.16) pm = |J j ^+(|)^+(^')/ (2 >^(r)^(0 d*d£'. (64.24) In particular, to a physical quantity /(£) that is simply a function of £, there corresponds the operator (64.23), which in this case can be written in the form f §(£ — 1 ) conventionally denotes the product 8(x-x )8(y-y Q )8(z-z )8 a §65 Second quantisation. The case of Fermi statistics 227 Hence it is clear that ^ + (£)^(£) d| is the operator of the number of particles in the range d|. When expressed by means of the operators ^F, ^ + , the operator H takes the form ti = j {(^ 2 /2m)VT+(0V^(l)+ U^)^ + (0^(i)} <*£+ + ij-J xr+(^^+(f)t/< 2 ^,f )^(f )"¥($) d^df + (64.25) We have used here the expression (64.18) for i? (1) , and have integrated by parts (with respect to the co-ordinates) the term containing the Laplacian. We can clarify the formula (64.25) by noticing the following point. Sup- pose that we have a system of particles, each of which is described (at a given instant) by the same wave function «/»(£), which we suppose normalised so that J \tjj\ 2 d£ = N. Then it is immediately evident that, if we replace the operator ty in the expression (64.25) by the function ifj, this expression becomes the mean energy of the system in the state considered. This gives the following rule for deriving the Hamiltonian in the second quantisation formalism. The expression for the mean energy is written in terms of the wave function of an individual particle (normalised as stated above), and this function is then replaced by the operator T, the Hermitian conjugate operators ^ + being written to the left of the operators T. If the system consists of bosons of different kinds, operators a, d + or ty, y i' + must be defined in the second quantisation method for each kind of particle. Operators pertaining to particles of different kinds of course commute. §65. Second quantisation. The case of Fermi statistics The basic theory of the method of second quantisation remains wholly unchanged for systems of identical fermions, but the actual formulae for the matrix elements of quantities and for the operators di are naturally different. The wave function Tjvat s ... now has the form (61.2): Vnn 1 2 VNl Mfor) 0,_(*l) 0. (& - *»-(**) (65.1) Because of the antisymmetry of this function, the question of its sign arises first of all. This question did not arise in the case of Bose statistics, since, 228 Identity of Particles §65 because of the symmetry of the wave function, its sign, once chosen, was preserved under all permutations of the particles. In order to make definite the sign of the function (65.1), we shall agree to choose it as follows. We number successively, once and for all, all the states fa. We then complete the rows of the determinant (65.1) so that always Px<p2<P 3 <-<pN, (65.2) whilst in the successive columns we have functions of the different variables in the order f lf £ 2 , ... , g N . No two of the numbers p^p 2 , ... can be equal, since otherwise the determinant would vanish. In other words, the occupa- tion numbers N t can take only the values and 1. Let us again consider an operator of the form (64.2), F a) = S/ (1) tt . As in §64, its matrix elements will be non-zero only for transitions where all the occupation numbers remain unchanged and for those where one occupation number (A^) is diminished by unity (becoming zero instead of one) and an- other (N k ) is increased by unity (becoming one instead of zero). We easily find that, for i < k, (F ll) Y<°* =/ t V-l)** w -*- 1 \ (65.3) V*; where by t , l t we signify N ( = 0, Nt = 1 and the symbol 2(&, /) denotes the sum of the occupation numbers of all states from the Mi to the /thrj- S (*,/)= S N». For the diagonal elements we obtain our previous formula (64.5) : FW = ?/w,AT,. (65.4) In order to represent the operator ^ (1) in the form (64.14), the operators d t must be defined as matrices whose elements are («*)£ = (^^(-l)* 1 -*-". (65.5) On multiplying these matrices, we find, for k > i, (ai+a^hh = («, + ) 1 A(a ft ) * * = (-1) I(1 « *'- 1) (- l)3i.*-iHS(*fi.*-i) °i 1 fc °i°k Vfc or (aSa^ = (- l)**w. *-». (65.6) If/ = k, the matrix of ^ + a i is diagonal, and its elements are unity for N t = 1, and zero for N t = ; this can be written _^^___ di+di = N t . (65.7) t For i > k the exponent in (65.3) becomes S (k + 1, i — 1). The sum must be taken as zero when i = k + 1. §65 Second quantisation. The case of Fermi statistics 229 On substituting these expressions in (64.14), we in fact obtain (65.3), (65.4). Multiplying a/*", d k in the opposite order, we have (a^+yi^ = (a k yi^(ai+yiH == (_l)i(i.^-i)+i+i(*fi.A-i)+^a.*-i)+i, or (ajflt+ffi* = -(- l) I(i+1 ' & - 1) . (65.8) Comparing (65.8) with (65.6), we see that these quantities have opposite signs, i.e. we can write di + d k +d k di + = (i # k). For the diagonal matrix a i a/ f , we find &.&* = i-N t . (65.9) Adding this to (65.7), we obtain didi + +di + di = 1. Both the above equations can be written in the form didif+dk+di = S ik . (65.10) On carrying out similar calculations, we find for the products d i d k the re- lations did k +d k di = 0, (65.11) and in particular d^ = 0. Thus we see that the operators d i and d k (or d k + ) for i # k anticommute, whereas in the case of Bose statistics they commuted with one another. This difference is perfectly natural. In the case of Bose statistics, the operators di and d k were completely independent; each of the operators d t acted only on a single variable N it and the result of this action did not depend on the values of the other occupation numbers. In the case of Fermi statistics, however, the result of the action of the operator d t depends not only on the number N t itself, but also on the occupation numbers of all the preceding states, as we see from the definition (65.5). Hence the action of the various operators d it d k cannot be considered independent. The properties of the operators d if d t + having been thus defined, all the remaining formulae (64.14)-(64.19) remain valid. The formulae (64.23)- (64.25), which express the operators of physical quantities in terms of the operators x F(f), 1 F + (£) defined by (64.20), also hold good. The commutation rules (64.21), (64.22), however, are now obviously replaced by 230 Identity of Particles §65 If the system consists of particles of different kinds, second quantisation operators must be defined for each kind of particle (as already mentioned at the end of §64). Operators belonging to bosons and fermions commute; those belonging to different fermions may formally be regarded as either commutative or anticommutative within the limits of non-relativistic theory. On either assumption the results obtained by means of the second quantisa- tion method are the same. However, with a view to later applications in the relativistic theory, which allows different particles to be transformed into one another, we should assume that the creation and annihilation operators for different fermions anticommute. This becomes evident if we regard as "different" particles two different "internal" states of a single complex particle. CHAPTER X THE ATOM §66. Atomic energy levels In the non-relativistic approximation, the stationary states of the atom are determined by Schrodinger's equation for the system of electrons, which move in the Coulomb field of the nucleus and interact electrically with one another; the spin operators of the electrons do not appear in this equation. As we know, for a system of particles in a centrally symmetric external field the total orbital angular momentum L and the parity of the state are conserved. Hence each stationary state of the atom will be characterised by a definite value of the orbital angular momentum L and by its parity. Moreover, the co-ordinate wave functions of the stationary states of a system of identical particles have a certain permutational symmetry. We have seen in §63 that, for a system of electrons, a definite value of the total spin of the system cor- responds to each type of permutational symmetry (i.e. to each Young dia- gram). Hence every stationary state of the atom is characterised also by the total spin S of the electrons. The converse, however, is of course not true; if L, S and the parity are given, the energy of the state is not uniquely determined. The energy level having given values of S and L is degenerate to a degree equal to the number of different possible directions in space of the vectors S and L. The degree of the degeneracy from the directions of L and S is re- spectively 2L + 1 and 2S + 1 . Consequently, the total degree of the degener- acy of a level with given L and S is equal to the product (2L + 1)(2S+1). In fact, however, there is always some relativistic electromagnetic inter- action of the electrons, which depends on their spins. It has the result that the energy of the atom depends not only on the absolute values of the orbital angular momentum and spin vectors, but also on their relative positions. Strictly speaking, when the relativistic terms in the Hamiltonian operator are taken into account, it no longer commutes with the operators L and §, i.e. the orbital angular momentum and the spin are not separately conserved. Only the total angular momentum J = L+S is conserved. The conservation of the total angular momentum is an exact law which follows at once from the isotropy of space relative to a closed system. For this reason the energy levels must be characterised by the values / of the total angular momentum. However, if the relativistic effects are comparatively small (as happens in many cases), they can be allowed for as a perturbation. Under the action of this perturbation, a level with given L and S, having (2L+l)(2S+l)-fold degeneracy, is "split" into a number of distinct (though close) levels, which differ in the value of the total angular momentum /. These levels are 231 232 The Atom §67 determined (in the first approximation) by the appropriate secular equation (§39), while their wave functions (in the zeroth approximation) are definite linear combinations of the wave functions of the initial degenerate level with the given L and S. In this approximation we can therefore, as before, regard the absolute values of the orbital angular momentum and spin (but not their directions) as being conserved, and characterise the levels by the values of L and S also. Thus, as a result of the relativistic effects, a level with given values of L and S is split into a number of levels with different values of /. This splitting is called the fine structure (or the multiplet splitting) of the level. As we know, J takes values from L+S to \L— S\ ; hence a level with given L and S is split into 25+1 (if L > S) or 2L + 1 (if L < S) distinct levels. Each of these is still degenerate with respect to the directions of the vector J; the degree of this degeneracy is 2/ + 1. It is easily verified that the sum of the numbers 2/ + 1 for all possible values of /is equal to (2L + l)(2S+l), as it should be. There is a generally accepted notation to denote the atomic energy levels (or, as they are called, the spectral terms of the atoms), similar to that used for the states of individual particles with definite values of the angular momentum (§32): states with different values of the total orbital angular momentum L are denoted by capital Latin letters, as follows : L = 0123456789 10... SPDFGHIKLMN ... Above and to the left of this letter is placed the number 2S+1, called the multiplicity of the term (though it must be borne in mind that this number gives the number of fine-structure components of the level only when L ^ S). Below and to the right of the letter is placed the value of the total angular momentum J. Thus the symbols 2 P 1/2 , 2 P 3/2 denote levels with L = 1, S = h J = \ and f . §67. Electron states in the atom An atom with more than one electron is a complex system of mutually interacting electrons moving in the field of the nucleus. For such a system we can, strictly speaking, consider only states of the system as a whole. Nevertheless, it is found that we can, with fair accuracy, introduce the idea of the states of each individual electron in the atom, as being the stationary states of the motion of each electron in some effective centrally symmetric field due to the nucleus and to all the other electrons. These fields are in general different for different electrons in the atom, and they must all be defined simultaneously, since each of them depends on the states of all the other electrons. Such a field is said to be self-consistent (see §69). Since the self-consistent field is centrally symmetric, each state of the elec- tron is characterised by a definite value of its orbital angular momentum /. The states of an individual electron with a given / are numbered (in order §67 Electron states in the atom 233 of increasing energy) by the principal quantum number n, which takes the values » = /+l,/+2, ... ; this choice of the order of numbering is made in accordance with what is usual for the hydrogen atom. However, it must be noticed that the sequence of levels of increasing energy for various / in com- plex atoms is in general different from that found in the hydrogen atom. In the latter, the energy is independent of /, so that the states with larger values of n always have higher energies. In complex atoms, on the other hand, the level with n= 5,1=0, for example, is found to lie below that with n = 4, 1=2 (this is discussed in more detail in §73). The states of individual electrons with different values of n and / are customarily denoted by a figure which gives the value of the principal quantum number, followed by a letter which gives the value of /:f thus \d denotes the state with n = 4, / = 2. A complete description of the atom demands that, besides the values of the total L, S and/, the states of all the electrons should also be enumerated. Thus the symbol 1* 2p 3 P denotes a state of the helium atom in which L = 1,S = 1,J = and the two electrons are in the Is and 2p states. If several electrons are in states with the same / and w, this is usually shown for brevity by means of an index: thus 3p 2 denotes two electrons in the 3p state. The distribution of the electrons in the atom among states with different / and n is called the electron configuration. For given values of n and /, the electron can have different values of the projections of the orbital angular momentum (m) and of the spin (a) on the #-axis. For a given /, the number m takes 2/+1 values; the number a is restricted to only two values, ±$. Hence there are altogether 2(2/+ 1) different states with the same n and /; these states are said to be equivalent. According to Pauli's principle there can be only one electron in each such state. Thus at most 2(2/+ 1) electrons in an atom can simultaneously have the same n and /. An assembly of electrons occupying all the states with the given n and / is called a closed shell of the type concerned. The difference in energy between atomic levels having different L and S but the same electron configuration:}: is due to the electrostatic interaction of the electrons. These energy differences are usually small, and several times less than the distances between the levels of different configurations. The following empirical principle (Hund's rule) is known concerning the relative position of levels with the same configuration but different L and S: The term with the greatest possible value of S (for the given electron con- figuration) and the greatest possible value of L (for this S) has the lowest energy. || f Another terminology often used is that in which electrons with principal quantum numbers n = 1, 2, 3, ... are said to belong to the K, L, M, . . . shells (see §74). J We here ignore the fine structure of each multiplet level. || The requirement that S should be as large as possible can be explained as follows. Let us consider, for example, a system of two electrons. Here we can have S = or S = 1 ; the spin 1 corresponds to an antisymmetrical co-ordinate wave function <f>(r lt r 2 ). For r x = r 2 , this function vanishes; in other words, in the state with 5=1 the probability of finding the two electrons close together is small. This means that their electrostatic repulsion is comparatively small, and hence the energy is less. Similarly, for a system of several electrons, the "most antisymmetrical" co-ordinate wave function corresponds to the greatest spin. 234 The Atom §67 We shall show how the possible atomic terms can be found for a given elec- tron configuration. If the electrons are not equivalent, the possible values of L and S are determined immediately from the rule for the addition of angular momenta. Thus, for instance, with the configurations np, n'p (n, ri being different) the total angular momentum L can take the values 2, 1,0, and the total spin S = 0, 1 ; combining these, we obtain the terms ^S, If we are concerned with equivalent electrons, however, restrictions im- posed by Pauli's principle make their appearance. Let us consider, for example, a configuration of three equivalent p electrons. For 1=1 (the p state), the projection m of the orbital angular momentum can take the values m = 1, 0, — 1, so that there are six possible states, with the following values of m and a: («)l,i (b)0,i (c)-l,i (a') 1,-1 (b') 0,-1 (0-l,-i. The three electrons can be one in each of any three of these states. As a result we obtain states of the atom with the following values of the projections M L = 2m, M s = £<r of the total orbital angular momentum and spin: {a+a'+b)2,\ ( a +a'+c)l,i (a+b+c) 0, f (a+b+b')l,i (a+b+c') 0,$ (a+b'+c)0,\ (a'+b+c)0,l The states with M L or M s negative need not be written out, since they give nothing different. The presence of a state with M L = 2, M s = \ shows that there must be a 2 D term, and to this term there must correspond one state (1, £) and one (0, \). Next, there remains one state with (1, Q, so that there must be a 2 P term; one of the states (0, £) corresponds to this. Finally, there remain the states (0, f ) and (0, £), corresponding to a 4 S term. Thus, for a configuration of three equivalent p electrons, the only possibilities are one term of each of the types 2 Z), 2 P, *S. Table 1 gives the possible terms for various configurations of equivalent p and d electrons. The figures below the letters of the terms show the num- ber of terms of the type concerned that exist for the given configuration, if this number is more than one. For the configuration with the greatest possible number of equivalent electrons (s 2 ,p 6 , d 10 , ...), the term is always X S. Like terms always correspond to configurations which differ in that one of them has as many electrons as the other lacks to form a closed shell. This is an evident result of the fact that the absence of an electron from the shell can be regarded as a "hole", whose state is defined by the same quantum numbers as the state of the missing electron. §67 Electron states in the atom Table 1 Possible terms for configurations of equivalent electrons 235 p,p s p*,p* P 3 *P 2 PZ> 3 P 4 5 d,d 3 d\d a d 3 ,d 7 1 SDG *PDFGH S PF *PF d^d* iSDFGI 2 2 2 3 PDFGH 2 2 5 D d 5 *SPDFGHI 3 2 2 *PDFG 6 S When Hund's rule is applied to determine the ground term of an atom from a known electron configuration, only the unfilled shell need be con- sidered, since the moments of electrons in closed shells cancel out. For example, let there be four d electrons outside the closed shells in an atom. The magnetic quantum number of the d electron can take five values: 0, ±1, ±2. Hence all four electrons can have the same spin component <j = |, and the maximum possible total spin is S = 2. We must then assign to the electrons different values of m so as to give the maximum value of M L = 2w, namely 2, 1, 0, — 1, M L = 2. This means that the maximum value of L for S = 2 is also 2, and the term is 5 D. PROBLEM Find the orbital wave functions of the possible states of a system of three equivalent p electrons. Solution. In the states 4 S the spins a of all the electrons are the same, and the values of m are therefore different. The wave function is given by a determinant of the form (61.2) composed of the functions ^o, 01, <A-i (where the suffix shows the value of w). For the 2 Z) term we consider the state with the maximum possible value Ml = 2. Two of the components m will be 1 and the other —1. Let electrons 2 and 3 have a = + \ and electron 1 have a = —\ (corresponding to total spin S = J). The orbital wave function having the required symmetry is = —-0i(l)[0o(2)^i(3)-0o(3)0i(2)] f V 2 the argument of each function being the number of the electron to which it refers. For the 2 P term we consider the state with Ml = 1 and the same values of the electron spin components as previously. This state can be obtained with two different sets of values of m, so that the orbital wave function is given by the linear combination j/j = aift-ui + bifjioo, 0-m = «Ai(l)[«A-i(2)0i(3)-«A-i(3)0i(2)], 0ioo = 0o(l)[0i(2)0o(3)-0i(3)0 o (2)]. 236 The Atom §68 To determine the coefficients, we use the relation Z+0 = (/+(!) + /+( 2 >+ /+(3))0 = 0, which must be satisfied by the wave function with Ml = M (see (27.8)). Using the matrix elements (27.12), we find that /+«Al = 0, l+ifj-i = -\/2i/j 0y t+ipo = \Z2ifjx, and so L+if>=^/2(a + b)ifj 011 = 0. Hence a + b = 0, and using also the normalisation condition, we have a = — b = \. The wave functions of states with Ml < L are obtained from those found above by apply- ing to them the operator £-. §68. Hydrogen-like energy levels The only atom for which Schrodinger's equation can be exactly solved is the simplest of all atoms, that of hydrogen. The energy levels of the hydro- gen atom, and of the ions He+, Li++, ... which each have only one electron, are given by Bohr's formula (36.10) mZ 2 e* 1 E = . _. (68.1) 2h%l + mJM) n 2 V ' Here Ze is the charge on the nucleus, M its mass, and m the mass of the elec- tron. We notice that the dependence on the mass of the nucleus is only very slight. The formula (68.1) does not take account of any relativistic effects. In this approximation there is an additional {accidental) degeneracy, peculiar to the hydrogen atom, of which we have already spoken in §36; for a given principal quantum number n, the energy is independent of the orbital angular momentum /. Other atoms have states whose properties recall those of hydrogen. We refer to highly excited states, in which one of the electrons has a large principal quantum number, and so is mostly at large distances from the nucleus. The motion of such an electron can be regarded, to a certain approximation, as motion in the Coulomb field of the rest of the atom, whose effective charge is unity. The values of the energy levels thus obtained are, however, too in- exact; it is necessary to apply to them a correction to take account of the devia- tion of the field from the pure Coulomb field at small distances. The nature of this correction is easily ascertained from the following considerations. Since the states with large quantum numbers are quasi-classical, the energy levels can be determined from Bohr and Sommerfeld's quantisation rule (48.6). The deviation from the Coulomb field at distances from the nucleus small compared with the "orbit radius" can be formally allowed for by an alteration in the boundary condition imposed on the wave function at t = 0. This brings about a change in the constant y in the quantisation condition for radial motion. Since this condition is otherwise unchanged, §69 The self-consistent field 237 we can conclude that we obtain for the energy levels an expression which differs from that for hydrogen in that the radial, that is, the principal, quantum number n is replaced by n + A/, where A* is some constant (known as Rydberg's correction) : m< £ 1 E = . (68.2) 1W- («+A,) 2 Rydberg's correction is (by definition) independent of n, but it is of course a function of the azimuthal quantum number / of the excited electron (which we add as a suffix to A), and of the angular momenta L and S of the whole atom. For given L and S, A* decreases rapidly as / increases. The greater /, the less time the electron spends near the nucleus, and hence the energy levels must approach more and more closely those of hydrogen as / increases.f §69. The self-consistent field Schrodinger's equation for atoms containing more than one electron can- not be directly solved in practice, even by numerical methods. Approximate methods of calculating the energies and wave functions of the stationary states of the atoms are therefore important. The most important of these methods is what is called the self-consistent field method. The idea of this method con- sists in regarding each electron in the atom as being in motion in the "self- consistent field" due to the nucleus together with all the other electrons. As an example, let us consider the helium atom, restricting ourselves to those terms in which both the electrons are in s states (with or without the same n) ; the states of the whole atom will then be S states also. Let ^(ri) and »/f 2 (r 2 ) be the wave functions of the electrons; in the s states they are functions only of the distances r v r 2 of the electrons from the nuclei. The wave function ^(r l5 r 2 ) of the atom as a whole is a symmetrised 4> = H^H^+H^H^) (69-1) or antisymmetrised = Ur^UH)-Mr*)Uri) (69 2) product of the two functions, according as we are concerned with states of t As an illustration, we may give the experimental values of Rydberg's correction for the highly excited states of the helium atom. The total spin of this atom can have the values S = and 1, while the total orbital angular momentum L is, in the states considered, the same as the angular momentum / of the excited electron (the other electron being in the state Is). Rydberg's corrections are for S = 0: A = -0-140, Ai = +0-012, A 2 = -00022; for S = 1 : A = -0-296, Ai = -0068, A 2 = -00029. 238 The Atom §69 total spinf S =. or S = 1 . We shall consider the second of these. The functions i/f x and i{j 2 can then be regarded as orthogonal. J Let us try to determine the function of the form (69.2) which is the best approximation to the true wave function of the atom. To do so, it is natural to start from the variational principle, allowing only functions of the form (69.2) to be considered; this method was proposed by V. A. Fok (1930). As we know, Schrodinger's equation can be obtained from the variational principle I i/t*lJipdV 1 dV 2 = minimum, with the additional condition (the integration is extended over the co-ordinates of both electrons in the helium atom). The variation gives the equation jj S<A*(#-£)«£ dV l( iV 2 = 0, (69.3) and hence, with an arbitrary variation of the wave function ip, we obtain the usual Schrodinger's equation. In the self-consistent field method, the expression (69.2) for ip is substituted in (69.3), and the variation is effected with respect to the functions ^ and ifj 2 separately. In other words, we seek an extremum of the integral with respect to functions iff of the form (69.2) ; as a result we obtain, of course, an inexact eigenvalue of the energy and an inexact wave function, but the best of the functions that can be represented in this form. The Hamiltonian for the helium atom is of the form|| 8 =#H-#,+ 1/ru, #x = -*Ai-2/r lf (69.4) where r 12 is the distance between the electrons. Substituting (69.2) in (69.3), carrying out the variation, and equating to zero the coefficients of 8^ and 8ip 2 in the integrand, we easily obtain the following equations: [lA+2/r+^-^ 22 -G 22 (r)]e/- 1 (r)+[H 12 +G 12 (r)]^ 2 (r) = 0, [*A+2/r+E-H 11 -G u (r)]^(r)+[ff 11 +G 11 (r)M 1 (r) =0, where * GaM = J «Aa(r 2 )<A b (r 2 ) dV 2 /r 12 , H ab = j 0.[-4A-2/r]& dV (a, b = 1, 2). (69.6) f The states of the helium atom with S = are usually called parahelium states, and those with S = 1 orthohelium states. J The wave functions tfi lt ip 2 , •■• of the various states of the electron which are obtained by the self- consistent field method are not in general orthogonal, since they are solutions of different equations, not of the same equation. In (69.2), however, without altering the function tp of the whole atom, we can replace tp 2 by tp 2 ' = 2 + constant X0 X ; by an appropriate choice of the constant, we can always ensure that ipj and tfi 2 ' are orthogonal. II In this section (including the Problems) we use atomic units (see the first footnote to §36). §69 The self-consistent field 239 These are the final equations resulting from the self-consistent field method; they can, of course, be solved only numerically, f The equations are similarly derived in more complex cases. The wave function of the atom to be substituted in the integral in the variational principle is in the form of a linear combination of products of the wave func- tions of the individual electrons. This combination must be so chosen that, firstly, its permutational symmetry corresponds to the total spin S of the state of the atom considered and, secondly, it corresponds to the given value of the total orbital angular momentum L of the atom. By using, in the variational principle, the wave function having the neces- sary permutational symmetry, we automatically take account of the exchange interaction of the electrons in the atom. Simpler equations (though leading to less accurate results) are obtained if we neglect the exchange interaction and also the dependence on L of the energy of the atom for a given electron configuration (D. R. Hartree 1928). As an example, let us again consider the helium atom; we can then write the equations for the wave functions of the electrons immediately in the form of ordinary Schrodinger's equations : [I Aa+E a - V a (r a )]Ur a ) =0 (a = 1, 2), (69.7) where V a is the potential energy of one electron moving in the field of the nucleus and in that of the distributed charge of the other electron: V 1 (r 1 ) = -2M- j (l!r 12 W(r 2 ) dV z , (69.8) and similarly for V % . In order to find the energy E of the whole atom, we must notice that, in the sum E x +E 2 , the electrostatic interaction between the two electrons is counted twice, since it appears in the potential energy Fifa) of the first electron and in that — V 2 (r 2 ) — of the second. Hence E is obtained from the sum E x -\-E^ by subtracting once the mean energy of this interaction; that is, E = E x +E 2 - Jj (l/r 12 W(r x ),A 2 2 (r 2 ) dV.dV,. (69.9) To refine the results obtained by this simplified method, the exchange interaction and the dependence of the energy on L can afterwards be taken into account as perturbations. PROBLEMS Problem 1. Determine approximately the energy of the ground level of the helium atom and helium-like ions (a nucleus of charge Z and two electrons), regarding the interaction between the electrons as a perturbation. Solution. In the ground state of the ion, both electrons are in s states. The unperturbed t A comparison of the energy levels of light atoms, calculated by the self-consistent field method, with spectroscopic data enables us to estimate the accuracy of the method at about 5 per cent. 240 The Atom §69 value of the energy is twice the ground level of a hydrogen-like ion (because of the two electrons) : EP» = 2{-\Z*)=-Z*. The correction in the first approximation is given by the mean value of the electron inter- action energy in a state with wave function Z3 77 (the product of two hydrogen functions with / = 0). The integral =/H EW= | 102 dPWa ri2 is most simply calculated as oo r 2 EP> = 2 J d V 2 . P2— fpid V h d Fi = 4 7 rr 1 2dr 1 , dV 2 = 477r 2 2 dr2, the energy of the charge distribution p 2 = I fa I 2 in the field of the spherically symmetric distribution pi = | fa |2; the integrand with dF 2 is the energy of the charge pn(r 2 ) in the field of the sphere n <r 2 , and the factor 2 takes account of the contribution from configurations in which n > r%. Thus we find E^ = 5Z/8, and finally E=Ef® +#!>= -Z2+|Z. For the helium atom (Z = 2) this gives —J? = 11/4 = 2-75; the actual value of the ground- state energy of this atom is — E = 2-90 atomic units = 78-9 eV. Problem 2. The same as Problem 1, but using the variational principle, approximating the wave function by a product of two hydrogen functions with some effective nuclear charge. Solution. We calculate the integral IS ifJlfdVidVz, #=-&Ai+A 2 ) +— n r 2 r 12 with the function tji given by (1), Problem 1 , but with Zett instead of Z. The integral of fa/m is calculated in Problem 1 ; the integral of if> Ai "A can be reduced to that of ^ 2 /n, since, by Schrodinger's equation, (-*Ai )<Al = -iZen^L The result is jjifsff l / J dV 1 dV 2 =Z efl 2-2ZZe fl + fZ efl . This expression as a function of Zett has a minimum at Zett = Z—&. The corresponding value of the energy is For the helium atom this gives — E = 2-85. It may be noted that the wave function (1) with the above value of Zett is in fact the best not only of all functions of the form (1) but of all functions which depend only on the sum r\ + r 2 . §70 The Thomas-Fermi equation 241 §70. The Thomas-Fermi equation Numerical calculations of the charge distribution and field in the atom by the self-consistent field method are extremely cumbersome, especially for complex atoms. For these, however, there is another approximate method, whose value lies in its simplicity; its results are admittedly much less accurate than those of the self-consistent field method. The basis of this method (E. Fermi, and L. Thomas, 1927) is the fact that, in complex atoms with a large number of electrons, the majority of the elec- trons have comparatively large principal quantum numbers. In these condi- tions the quasi-classical approximation is applicable. Hence we can apply the concept of "cells in phase space" (§48) to the states of the individual electrons. The volume of phase space corresponding to electrons which have momenta less than p and are in the volume element dV of physical space is iirp z dV. The number of "cells", i.e. possible states, corresponding to this volume isf 4ttP 3 dVjZilrtf, and in these states there cannot at any one time be more than 4it/> 3 p z 2-J—dV = ^—dV 3(2t7) 3 3?t 2 electrons (two electrons, with opposite spins, in each "cell"). In the normal state of the atom, the electrons in each volume element dV must occupy (in phase space) the cells corresponding to momenta from zero up to some maxi- mum value p . Then the kinetic energy of the electrons will have its smallest possible value at every point. If we write the number of electrons in the volume dV as ndV (where n is the number density of electrons), we can say that the maximum value Po of the momenta of the electrons at every point is related to n by P*jZtt* = n. The greatest value of the kinetic energy of an electron at a point where the electron density is n is therefore W=\{Wnyi*. (70.1) Next, let ${r) be the electrostatic potential, which we suppose zero at infinity. The total energy of the electron is ip 2 —<f>. It is evident that the total energy of each electron must be negative, since otherwise the electron moves off to infinity. We denote the maximum value of the total energy of the electron at each point by — <£ , where <£ is a positive constant ; if this quan- tity were not constant, the electrons would move from points with smaller <j> to those with greater <f> . Thus we can write W = <t>~K (70.2) t In this section we use atomic units. 242 The Atom §70 Equating the expressions (70.1) and (70.2), we obtain n = [2(^ )] 3 / 2 /37r 2 , (70.3) a relation between the electron density and the potential at every point in the atom. For <j>=(j> Q the density n vanishes ; n must clearly be put equal to zero also in the whole of the region where <f> < <j> , and where the relation (70.2) would give a negative maximum kinetic energy. Thus the equation <f> = <f> deter- mines the boundary of the atom. There is, however, no field outside a centrally symmetric system of charges whose total charge is zero. Hence we must have <f> = at the boundary of a neutral atom. It follows from this that, for a neutral atom, the constant <£ must be put equal to zero. On the other hand, <f> is not zero for an ion. Below we shall consider a neutral atom, putting accordingly <j> Q = 0. According to Poisson's electrostatic equation, we have A<f> = 4irn; substitut- ing (70.3) in this, we obtain the fundamental equation of the Thomas-Fermi method : A«£ = (8V2/377-)^/ 2 . (70.4) The field distribution in the normal state of the atom is determined by the centrally symmetric solution of this equation that satisfies the following boundary conditions: for r -> the field must become the Coulomb field of the nucleus, i.e. (f>r -> Z, while for r -> oo we must have^r -» 0. Introducing here, in place of the variable r, a new variable x according to the definitions r = xbZ-y*, b = Kf^) 2/3 = 0-885, (70.5) and, in place of (f>, a new unknown function x by f Z /rZ l l*\ ZW x(x) <f>( r )=-x(-ir)=-T- — • < 70 - 6 ) r \ b J b x we obtain the equation x i>2 tf x jdx 2 = x s/2 , (70.7) with the boundary conditions x = 1 for ^ = and x = for x = oo. This equation contains no parameters, and thus defines a universal function x(x). Table 2 gives values of this function obtained by numerical integration of equation (70.7). The function x(x) decreases monotonically, and vanishes only at infinity 4 In other words, the atom has no boundaries in the Thomas- Fermi model, and formally extends to infinity. t In ordinary units, <f>(r) = {Ze\r)x{rZV*me*\W&5W-). % The equation (70.7) has the exact solution \( x ) — 144x~ 3 , which vanishes at infinity but does not satisfy the boundary condition at x = 0. It could be used as an asymptotic expression for the function X(x) for large x. However, this expression gives fairly exact values only for very large x, whilst the Thomas-Fermi equation becomes inapplicable at large distances (see below). §70 The Thomas-Fermi equation Table 2 Values of the function x(x) 243 X X(x) X X(x) X X(x) 0-00 1-000 1-4 0-333 6 0-0594 0-02 0-972 1-6 0-298 7 0-0461 0-04 0-947 1-8 0-268 8 0-0366 0-06 0-924 2-0 0-243 9 0-0296 0-08 0-902 2-2 0-221 10 0-0243 0-10 0-882 2-4 0-202 11 0-0202 0-2 0-793 2-6 0-185 12 0-0171 0-3 0-721 2-8 0-170 13 0-0145 0-4 0-660 3-0 0-157 14 0-0125 0-5 0-607 3-2 0-145 15 0-0108 0-6 0-561 3-4 0-134 20 0-0058 0-7 0-521 3-6 0-125 25 0-0035 0-8 0-485 3-8 0-116 30 0-0023 0-9 0-453 4-0 0-108 40 0-0011 1-0 0-424 4-5 0-0919 50 0-00063 1-2 0-374 5-0 0-0788 60 0-00039 The value of the derivative x '(x) for x = is x'(°) = —1-59. Hence, as x -> 0, the function x(%) is of the form x = 1 — 159*, and accordingly the potential <f>(r) is <f>(r) ^ Z/r-l-80Z 4 / 3 . (70.8) The first term is the potential of the field of the nucleus, while the second ( — \%0me z Z m jh 2 ' in ordinary units) is the potential at the origin due to the electrons. Substituting (70.6) in (70.3), we find for the electron density an expression of the form n = Z*f{rZ™\b\ f(x) = (32/9tt%/*)3/2. (70.9) We see that, in the Thomas- Fermi model, the charge density distribution in different atoms is similar, with Z~ 1/3 as the characteristic length (in ordinary units h 2 /me 2 Z 1/3 , i.e. the Bohr radius divided by Z 1/3 ). If we measure distances in atomic units, the distances at which the electron density has its maximum value are the same for all Z. Hence we can say that the majority of the electrons in an atom of atomic number Z are at distances from the nucleus of the order of Z~ 1/3 . A numerical calculation shows that half the total electron charge in an atom lies inside a sphere of radius 1-33 Z~ 1/3 . Similar considerations show that the mean velocity of the electrons in the 9 244 The Atom §70 atom (taken, as an order of magnitude, as the square root of the energy) is of the order of Z 2/3 . The Thomas- Fermi equation becomes inapplicable both at very small and at very large distances from the nucleus. Its range of applicability for small r is restricted by the inequality (49.12); at smaller distances the quasi-classical approximation becomes invalid in the Coulomb field of the nucleus. Putting in (49.12) a = Z, We find 1/Z as the lower limit of distance. The quasi- classical approximation becomes invalid for large r also in a complex atom. In fact, it is easy to see that, for r ~ 1, the de Broglie wavelength of the elec- tron becomes of the same order of magnitude as the distance itself, so that the quasi-classical condition is undoubtedly violated. This can be seen by esti- mating the values of the terms in equations (70.2) and (70.4); indeed, the result is obvious without calculation, since equation (70.4) does not involve Z. Thus the applicability of the Thomas-Fermi equation is limited to dis- tances large compared with \\Z and small compared with unity. In complex atoms, however, the majority of the electrons in fact lie in this region. This means that the "outer boundary" of the atom in the Thomas-Fermi model is at r~ 1, i.e. the dimensions of the atom do not depend on Z. The energy of the outer electrons, i.e. the ionisation potential of the atom, is likewise independent of Z.f By means of the Thomas-Fermi method we can calculate the total ionisa- tion energy E y i.e. the energy needed to remove all the electrons from the neutral atom. To do this, we must calculate the electrostatic energy of the Thomas-Fermi distribution for the charges in the atom; the required total energy is half this electrostatic energy, since the mean kinetic energy in a system of particles interacting in accordance with Coulomb's law is (by the virial theorem) minus half the mean potential energy. The dependence of E on Z can be determined a priori from simple considerations: the electrostatic energy of Z electrons at a mean distance Z -1/3 from a nucleus of charge Z, and moving in its field, is proportional to Z Z\Z~ VZ = Z 7/3 . A numerical calculation gives the result E = 20-8Z 7/3 eV. The dependence on Z is in good agreement with the experimental data, though the empirical value of the coefficient is close to 16. We have already mentioned that positive (non-zero) values of the constant cf> correspond to ionised atoms. If we define the function x by <f> —<f> = Z%\r y we obtain the same equation (70.7) for x as previously. We must now, how- ever, consider only solutions which vanish not at infinity as for the neutral atom, but for finite values x of x. Such solutions exist for any x . At the point x = x 0f the charge density vanishes together with x, but the potential remains finite. The value of x is related to the degree of ionisation in the following manner. The total charge inside a sphere of radius r is, by Gauss's t This model does not, of course, show the periodic dependence of the dimensions and ionisation potential of the atom on Z, which appears in the periodic system of the elements. Moreover, experi- mental data indicate the existence of a slight but steady increase in dimensions and decrease in the ionisation potential as Z increases. §70 The Thomas-Fermi equation 245 theorem, —r 2 d(f>jdr = Z[x(x)— xx'(x)]. The total charge z on the ion is obtained by putting x = x in this ; since x( x o) = 0» we have z = -Zx x'(x ). (70.10) The thick line in Fig. 23 shows the curve of x(x) for a neutral atom; below it are two curves for ions of different degrees of ionisation. The quantity z\Z is shown graphically by the length of the segment intercepted on the axis of ordinates by the tangent to the curve at x = x . Fig. 23 Equation (70.7) also has solutions which are nowhere zero; these diverge at infinity. They can be regarded as corresponding to negative values of the constant (f> . Fig. 23 also shows two such curves of x(x) ; they lie above the curve for the neutral atom. At the point x = x lt where X(*i)-*i*'(*i) = 0, (70.11) the total charge inside the sphere x < x x is zero (graphically, this point is evidently the one where the tangent to the curve passes through the origin). If we cut off the curve at this point, we can say that it defines x(x) for a neutral atom at whose boundary the charge density remains non-zero. Physically, this corresponds to a "compressed" atom confined to some given finite volume.f The Thomas- Fermi equation does not take account of the exchange inter- action between electrons. The effects which this involves are of the next order of magnitude with respect to Z~ 2/3 . Hence an allowance for the ex- change interaction in the Thomas-Fermi method requires a simultaneous consideration of both these effects and others of the same order of magnitude. J t This approach may be useful in studying the equation of state of highly compressed matter. t This has been done by A. S. Kompaneets and E. S. Pavlovskii (Soviet Physics JETP 4 328 1957) and by D. Kirzhnits (ibid. 5, 64, 1338, 1957). 246 The Atom §71 PROBLEM Find the relation between the energy of the electrostatic interaction between electrons and that of their interaction with the nucleus in a neutral atom, using the Thomas- Fermi model. Solution. The potential 4> e of the field due to the electrons is found by subtracting the potential Zjr of the nucleus front the total potential 4>. The energy of the interaction between the electrons is therefore = $z(-?-dV-$[<l>ndV -**Jr ,r - L T-J : (3tt2)2/3 (where <f> has been expressed in terms of n by means of (70.3)). The energy U e n of the interaction between the electrons and the nucleus and their kinetic energy T are therefore U en =-ZJ^dV, (3tt2)2/3 = 3- — » 5 /3dF. 10 ■/■ Comparing these expressions with the previous equation, we find Uee = — $Uen e-** According to the virial theoremf, for a system of particles interacting according to Coulomb's law we have IT = — U = —Uen — Uee. Thus finally §71. Wave functions of the outer electrons near the nucleus We have seen, on the basis of the Thomas- Fermi model, that the outer electrons in complex atoms (Z large) are mainly at distances r ~ 1 from the nucleus. $ A number of properties of atoms, however, depend significantly on the electron density near the nucleus; such properties will be considered in §§72 and 120. To determine the order of magnitude of this density we may examine the variation of the wave function ?/»(r) of the electron in the atom when r varies from large (r ~ 1) to small distances. t See Mechanics, §10. j In this section we use atomic units. §72 Fine structure of atomic levels 247 In the region r ~ 1, the field of the nucleus is screened by the remaining electrons, so that the potential energy U{r) ~ 1/r ~ 1. The energy of the electron level in this field E ~ 1. At distances of the order of the Bohr radius in the field of a charge Z, r ~ 1/Z, the field of the nucleus may be regarded as unscreened, and U = — Zjr. In the transitional region, 1/Z < r < 1, the potential energy j t7 j is large compared with the electron energy E, and the condition d 1 1(1) dr\pj pj dr ,/\U\ holds (where/> is the momentum), so that the motion of the electron is quasi- classical. The spherically symmetrical quasi-classical wave function is 1 1 1 U( r )| f or _ <^1, (71.1) the order of magnitude of the coefficient (~ 1) being determined by the con- dition «/r ~ 1 for "joining" to the wave function for r ~ 1. Applying the expression (71.1) in order of magnitude for r ~ 1/Z (sub- stituting U = — Z/r), we obtain the required value of the wave function near the nucleus :f 4l\\Z)~y/Z. (71.2) In accordance with the general properties of wave functions in a central field (§32), when the distance decreases further ift(r) either remains constant in order of magnitude (for an s electron) or begins to decrease (for 7^0). The probability of finding the electron in the region r < 1/Z is w~\iP\W~l/Z2. (71.3) The formulae (71.2) and (71.3) of course determine only the systematic variation with increasing Z, and do not take into account non-systematic variations from one element to the next. §72. Fine structure of atomic levels A detailed study of relativistic interactions will be given in Volume 4, but some properties of such interactions may be mentioned here. It is found that the relativistic terms in the Hamiltonian of an atom fall into two classes. One of these contains terms linear with respect to the spin operators of the electrons, while the other includes quadratic terms. The former correspond to the interaction between the orbital motion of the electrons and their spin (this interaction is called spin-orbit interaction), while the latter correspond to the interaction between the spins of the electrons (spin-spin interaction). t To determine the coefficient in this formula (when the wave function is known in the region r ~ 1), we should have to use the expression (36.25) in the range r ;S 1/Z. 248 The Atom §72 Both interactions are of the same order (the second) with respect to vjc y the ratio of the velocity of the electrons to that of light ; in practice, the spin- orbit interaction considerably exceeds the spin-spin interaction in heavy atoms. This is because the spin-orbit interaction increases rapidly with the atomic number, whereas the spin-spin interaction is essentially inde- pendent of Z (see below). The spin-orbit interaction operator is of the form ^ = SA s .s 8 (72.1) (the summation being over all the electrons in the atom), where s a are the spin operators of the electrons, and A a are some "orbital" operators, i.e. opera- tors acting on functions of the co-ordinates. In the self- consistent field approximation the operators A a are proportional to the operators \ a of the orbital angular momentum of the electrons, and V s i can then be written in the form V s i = 2a a l a . s a . (72.2) The coefficients in the sum are given in terms of the potential energy U(r) of the electron in the self-consistent field by h 2 dU(r a ) 2m 2 c 2 r a dr a (72.3) Since U < and | U(r)\ decreases away from the nucleus, all the <x a > 0. Regarding the interaction (72.1) as a perturbation, we should, in order to calculate the energy, average it with respect to the unperturbed state. The main contribution to the energy is given by distances close to the nucleus, of the order of the Bohr radius ( ~ h 2 /Zme 2 ) for a nucleus with charge Ze. In this region the field of the nucleus is almost unscreened and the potential energy is £/(r)~Ze 2 /r~Z 2 »*e 4 //* 2 , so that a.~h 2 Ulm 2 c 2 r 2 ~Z\e 2 \hcfm<A\h 2 . The mean value of a is obtained by multiplying by the probability w of finding the electron near the nucleus. According to (71.3), w~Z~ 2 , so that we have finally that the energy of the spin-orbit interaction of the electron is given by / Ze 2 \ 2 me 4 \ he J h 2 i.e. differs from the fundamental energy of the outer electrons in the atom (~me*/h 2 ) only by the factor (Ze 2 /hc) 2 . This factor increases rapidly with the atomic number, and reaches values of the order of unity in heavy atoms. §72 Fine structure of atomic levels 249 The actual averaging of the operator (72.1) is done in two steps. First of all, we average over electron states with given absolute values L and S of the total orbital angular momentum and spin, but not with given directions of these. After this averaging V' sl is still, of course, an operator, which we de- note by f Psl' From considerations of symmetry it is evident that the mean values of s a must be "directed" along S, which is the only spin "vector" characterising the atom as a whole (it must be recalled that, in the zeroth ap- proximation, the wave functions are products of a spin part and a co-ordinate part). Similarly, the mean values of l a must be "directed" along L. Thus the operator V~ 8L is of the form V 8 L=A%l h (72.4) where A is a constant characterising a given (unsplit) term, i.e. depending on S and L but not on the total angular momentum / of the atom. To calculate the energy of the splitting of a degenerate level (with given S and L), we must solve the secular equation formed from the matrix elements of the operator (72.4). In this case, however, we already know the correct functions in the zeroth approximation, in which the matrix of V$l is diagonal. These are the wave functions of states with definite values of the total angular momentum J. The averaging with respect to such a state involves replacing the operator S.L by its eigenvalues, which, according to the general formula (31.2), are L-S =IU(J+1)-L(L+1)-S(S+1)]. Since the values of L and S are the same for all the components of a multiplet, and we are interested only in their relative position, we can write the energy of the multiplet splitting in the form 44/C7+1). (72.5) The intervals between adjacent components (with numbers/ and /— 1) are consequently AEjj-! = AJ. (72.6) This formula gives what is called Lande's interval rule (1923). The constant A can be either positive or negative. For A > the lowest component of the multiplet level is the one with the smallest possible /, i.e. J = \L— S\ ; such multiplets are said to be normal. If A < 0, on the other hand, the lowest level of the multiplet is that with J = L+S; these multi- plets are said to be inverted. t This averaging signifies essentially the construction of a matrix with elements (nM'iM's Wn\ nMiMs) with all possible Mi, M'l and Ms, M's and diagonal with respect to all the other quantum numbers (the assembly of which we denote by n). 250 The Atom §72 It is easy to determine the sign of A for the normal states of atoms if the electron configuration is such that there is only one shell not completely filled. If this shell is not more than half filled, then according to Hund's rule (§67) all n electrons in it have parallel spins, so that the total spin has the greatest possible value, S = \n. Substituting in (72.2) s a = S/n and taking cc a (which is the same for all electrons in a given shell) outside the sum we obtain fitt = («/2S)8.L, i.e. A = a/2»S> 0. If the shell is more than half full, we first add and sub- tract in (72.2) the same sum taken over the unoccupied places or "holes" in the incomplete shell. Since, for a completely filled shell, we should have V s i = 0, the operator frsi is thereby represented as a sum V'si = -2 <x a l a .s a , taken only over the "holes", the total spin and orbital angular momentum of the atom being S = — 2 s a , L = — 2 l a . By the same method as pre- viously we therefore find A — — <x./2S, i.e. A < 0. From the above we have a simple rule which gives the value of / in the normal state of an atom with one incompletely filled shell. If this shell contains not more than half the greatest possible number of electrons for that shell, then/ = \ L-S\; if the shell is more than half full, J = L + S. As already mentioned, the spin-spin interaction, unlike the spin-orbit interaction, is essentially independent of Z. This is evident from the fact that it is a direct interaction between electrons and does not involve the field of the nucleus. For the averaged spin-spin interaction operator we should obtain, analog- ously to formula (72.4), an expression quadratic in S. The expressions S 2 and (S.L) 2 are quadratic in §. The former has eigenvalues independent of J, and therefore does not give any splitting of the term. Hence it can be omitted, and we can write P SS =:B(S.L)\ (72.7) where B is a constant. The eigenvalues of this operator contain terms inde- pendent of J, terms proportional to J(J+l), and finally a term proportional to J\J+ 1) 2 . The first of these do not give any splitting and hence are without interest; the second can be included in the expression (72.5), which simply means a change in the constant A. Finally, the last term gives an energy iB/ 2 C/+l) 2 . (72.8) The scheme for the construction of the atomic levels discussed in §§66-67 §72 Fine structure of atomic levels 251 is based on the supposition that the orbital angular momenta of the electrons combine to give the total orbital angular momentum L of the atom, and their spins to give the total spin S. As has already been mentioned, this supposi- tion is legitimate only when the relativistic effects are small; more exactly, the intervals in the fine structure must be small compared with the differences between levels with different L and S. This approximation is called the Russell- Saunders case, and we speak also of LS coupling. In practice, however, this approximation has a limited range of applica- bility. The levels of the light atoms are arranged in accordance with the LS model, but as the atomic number increases the relativistic interactions in the atom become stronger, and the Russell-Saunders approximation becomes inapplicable.f It must also be noticed that this approximation is, in parti- cular, inapplicable to highly excited levels, in which the atom contains an electron which is in a state with large «, and which is therefore mainly at large distances from the nucleus (see §68). The electrostatic interaction of this electron with the motion of the other electrons is comparatively weak, but the relativistic interaction in the rest of the atom is not diminished. In the opposite limiting case the relativistic interaction is large compared with the electrostatic (or, more precisely, compared with that part of it which governs the dependence of the energy on L and S). In this case we cannot speak of the orbital angular momentum and spin separately, since they are not conserved. The individual electrons are characterised by their total angular momenta/, which combine to give the total angular momentum/ of the atom. This scheme of arrangement of the atomic levels is called jj coupl- ing. In practice, this coupling is not found in the pure state, but various types of coupling intermediate between LS and jj are observed among the levels of very heavy atoms. J A peculiar type of coupling is observed in certain highly excited states. Here the rest of the atom may be in a Russell-Saunders state, i.e. may be characterised by the values of L and S, while its coupling with the highly excited electron is of the;)" type ; this is again due to the weakness of the elec- trostatic interaction for this electron. The fine structure of the energy levels of the hydrogen atom has certain characteristic properties. It will be calculated exactly in Volume 4, but here we shall only mention that, for a given principal quantum number n, the energy depends only on the total angular momentum j of the electron. Thus the degeneracy of the levels is not completely removed; to a level with given n and/ there correspond two states with orbital angular momenta I = j ± | (unless/ has the value « — -J, which is the greatest possible for a given n). t Nevertheless, it must be mentioned that, although the quantitative formulae which describe this type of coupling become inapplicable, the method of classifying levels according to this scheme may itself remain meaningful for heavier atoms, especially for the lowest states (including the normal state). % For further details regarding types of coupling and the quantitative aspect of the problem, see, for instance, E. U. Condon and G. H. Shortley, The Theory of Atomic Spectra, Cambridge University Press 1935 252 The Atom §73 Thus the level with n = 3 is split into three levels, of which the states S\I2, pi/2 correspond to one, ^3/2 and ^3/2 to another, and ^5/2 to the third. §73. The periodic system of D. I. Mendeleev The elucidation of the nature of the periodic variation of properties, ob- served in the series of elements when they are placed in order of increasing atomic number, requires an examination of the peculiarities in the successive completion of the electron shells of atoms. The theory of the periodic system is due to N. Bohr (1922). When we pass from one atom to the next, the charge is increased by unity and one electron is added to the envelope. At first sight we might expect the binding energy of each of the successively added electrons to vary monotonically as the atomic number increases. The actual variation, how- ever, is entirely different. In the normal state of the hydrogen atom there is only one electron, in the Is state. In the atom of the next element, helium, another Is electron is added; the binding energy of the Is electrons in the helium atom is, however, considerably greater than in the hydrogen atom. This is a natural conse- quence of the difference between the field in which the electron moves in the hydrogen atom and the field encountered by an electron added to the He + ion. At large distances these fields are approximately the same, but near the nucleus with charge Z = 2 the field of the He + ion is stronger than that of the hydrogen nucleus with Z = 1. In the lithium atom (Z = 3), the third electron enters the 2s state, since no more than two electrons can be in Is states at the same time. For a given Z the 2s level lies above the Is level; as the nuclear charge increases, both levels become lower. In the transition from Z = 2 to Z = 3, however, the former effect is predominant, and so the binding energy of the third electron in the lithium atom is con- siderably less than those of the electrons in the helium atom. Next, in the atoms from Be (Z = 4) to Ne (Z = 10), first one more 2s electron and then six 2p electrons are successively added. The binding energies of these electrons increase on the average, owing to the increasing nuclear charge. The next electron added, on going to the sodium atom (Z = 11), enters the 3s state, and the binding energy again diminishes markedly, since the effect of going to a higher shell predominates over that of the increase of the nuclear charge. This picture of the filling up of the electron envelope is characteristic of the whole sequence of elements. All the electron states can be divided into successively occupied groups such that, as the states of each group are occu- pied in a series of elements, the binding energy increases on the average, but when the states of the next group begin to be occupied the binding energy decreases noticeably. Fig. 24 shows those ionisation potentials of elements that are known from spectroscopic data; they give the binding energies of the electrons added as we pass from each element to the next. §73 The periodic system of D. I. Mendeleev 253 254 The Atom §73 The different states are distributed as follows into successively occupied groups : 1* 2 electrons \ 2s, Ip 8 3s , Zp 8 4s , 3 d, \p 18 Is, 4d t Sp 18 (73.1) 6s, 4/, Sd, dp 32 7*,6<f,5/,... / The first group is occupied in H and He; the occupation of the second and third groups corresponds to the first two (short) periods of the periodic system, containing 8 elements each. Next follow two long periods of 18 ele- ments each, and a long period containing the rare-earth elements and 32 elements in all. The final group of states is not completely occupied in the natural (and artificial transuranic) elements. To understand the variation of the properties of the elements as the states of each group are occupied, the following property of d and / states, which distinguishes them from s and/) states, is important. The curves of the effec- tive potential energy of the centrally symmetric field (composed of the electro- static field and the centrifugal field) for an electron in a heavy atom have a rapid and almost vertical drop to a deep minimum near the origin; they then begin to rise, and approach zero asymptotically. For s and p states, the rising parts of these curves are very close together. This means that the electron is at approximately the same distance from the nucleus in these states. The curves for the d states, and particularly for the / states, on the other hand, pass considerably further to the left; the "classically accessible" region which they delimit ends considerably closer in than that for the 5 and^ states with the same total electron energy. In other words, an electron in the d and/ states is mainly much closer to the nucleus than in the s and p states. Many properties of atoms (including the chemical properties of elements; see §81) depend principally on the outer regions of the electron envelopes. The above characteristic of the d and /states is very important in this connection. Thus, for instance, when the 4/ states are being filled (in the rare-earth ele- ments ; see below), the added electrons are located considerably closer to the nucleus than those in the states previously occupied. As a result, these electrons have practically no effect on the chemical properties, and all the rare-earth elements are chemically very similar. The elements containing complete d and / shells (or not containing these shells at ail) are called elements of the principal groups; those in which the filling up of these states is actually in progress are called elements of the inter- mediate groups. These groups of elements are conveniently considered separ- ately. §73 The periodic system of D. I. Mendeleev 255 Let us begin with the elements of the principal groups, helium have the following normal states: Hydrogen and iH:b 2 5 1/2 2 He:ls 21 »So (the number with the chemical symbol always signifies the atomic number). The electron configurations of the remaining elements of the principal groups are shown in Table 3. In each atom, the shells shown on the right of the table in the same line and above are completely filled. The electron configuration in the shells that are being filled is shown at the top, while the principal quantum number of the electrons in these states is shown by the figure on the left of the table in the same line. The normal states of the whole atom are shown at the bot- tom. Thus, the aluminium atom has the electron configuration Is 2 2s 2 2p* 3s 2 The values of L and S in the normal state of the atom can be determined (the electron configuration being known) by means of Hund's rule (§67), and the value of/ is determined by the rule given in §72. Table 3 Electron configurations of the atoms of elements in the principal groups 5 s* s*p s*p* 5 4 i>» *V s*p s s*p* n = 2 3Li 4 Be 5 B eC 7N 8 o 9F ioNe U* 3 iiNa 12Mg 13AI i 4 Si 15P ioS 17C1 iaAr 2s* 2p* 4 ioK 2oCa 3s 2 3p* 4 29C11 3oZn 3iGa 32Ge 33AS 34Se 3 5 Br 3eKr 3d 10 5 37 Rb ssSr 4s a Ap* 5 47Ag 48Cd 49I11 soSn 5iSb 52Te 53I 54Xe 4d 10 6 55CS 56Ba 5s* 5p* 6 79AU soHg 81TI 8 2 Pb 83Bi 84P0 85At 8eRn 4/ 1 * 5d l ° 7 87Fr ssRa 6s % 6p* *s 1/t 'So ■Pi/» 'Po 4 S./« 'P* *Pt/t l S The atoms of the inert gases (He, Ne, Ar, Kr, Xe, Rn) occupy a special position in the table: the filling up of one of the groups of states listed in (73.1) is completed in each of them. Their electron configurations have unusual stability (their ionisation potentials are the greatest in their respective series). This causes the chemical inertness of these elements. We see that the occupation of different states occurs very regularly in the series of elements of the principal groups : first the s states and then the p states are occupied for each principal quantum number n. The electron configurations of the ions of these elements are also regular (until electrons from the d and /shells are removed in the ionisation): each ion has the con- figuration corresponding to the preceding atom. Thus, the Mg+ ion has the configuration of the sodium atom, and the Mg++ ion that of neon. 256 The Atom §73 Let us now turn to the elements of the intermediate groups. The filling up of the 3d, 4d, and Sd shells takes place in groups of elements called respectively the iron group, the palladium group and the platinum group. Table 4 gives those electron configurations and terms of the atoms in these groups that are known from experimental spectroscopic data. As is seen from this table, the d shells are filled up with considerably less regularity than the s and p shells in the atoms of elements of the principal groups. Here a Table 4 Electron configurations of the atoms of elements in the iron, palladium and platinum groups Iron group 2lSc 22 Ti 23V 24Cr 25Mn 26Fe 27C0 28Ni Ar envelope + 3d As 2 2 D 3/2 3d 2 4s 2 *F 2 3d* 4s 2 *F 3/2 3d 6 4s 'S 3 3d* 4s* <s 5/2 3d 9 4s 2 8 #4 3d 7 4s 2 *F tlt 3d 6 4s 2 *F t Palladium group 39Y 4oZr 4lNb 42M0 43TC 44RU 4sRh 4ePd Kr envelope + 4d5s 2 2 D 3/2 4<f 2 5s 2 S F 2 4d*5s "A/2 4d 5 5s 'S 3 4rfs 5s 2 6 5 6 /2 4J' 5s 5 F 6 4d*5s *F 9/2 4d 10 'S Platinum group 57La Xe envelope + 5d6s s *D 3/2 7iLu 7 2 Hf 73Ta 74W 7sRe 7eOs nlr 7 8 Pt Xe envelope 1 +4/"+ ) 5 d 6s 2 2 D 3/2 5d* 6s 2 5d* 6s 2 4 F 3 /2 5d l 6s 2 6 D 5d 5 6s 2 e s 6/2 5d« 6s 2 5 £>4 5J7 6s 2 5d°6s *D 3 characteristic feature is the "competition" between the s and d states. It is seen in the fact that, instead of a regular sequence of configurations of the type dP s 2 with increasing/), configurations of the type d& +1 s or dv+ 2 are often found. Thus, in the iron group, the chromium atom has the configuration 3d 5 4s, and not 3d 4 4s 2 ; after nickel with 8 d electrons, there follows at once the copper atom with a completely filled d shell (and hence we place this §73 The periodic system of D. I. Mendeleev 257 element in the principal groups). This lack of regularity is observed in the terms of ions also: the electron configurations of the ions do not usually agree with those of the preceding atoms. For instance, the V+ ion has the configuration 3d* (and not 3d 2 4-s 2 like titanium) ; the Fe+ ion has 3d 6 4s (instead of 3d 5 4s 2 as in manganese). We may remark that all ions found naturally in crystals and solutions contain only d (not s or p) electrons in their incomplete shells. Thus iron is found in crystals or solutions only as the ions Fe++ and Fe +++ , whose configurations are 3d 6 and 3d 5 respectively. A similar situation occurs in the filling up of the 4/ shell ; this takes place in the series of elements known as the rare earths (Table 5).f The filling up of the 4/ shell also occurs in a slightly irregular manner characterised by the "competition" between 4/, Sd and 6s states. The last group of intermediate elements begins with actinium. In this group the 6d and 5/ shells are filled, similarly to what happens in the group of rare- earth elements (Table 6). To conclude this section, let us examine an interesting application of the Thomas-Fermi method. We have seen that the electrons in the p shell first appear in the fifth element (boron), the d electrons for Z = 21 (scandium), and the/electrons for Z = 58 (cerium). These values of Z can be predicted by the Thomas-Fermi method, as follows. An electron with orbital angular momentum / in a complex atom moves with an "effective potential energy" J of The first term is the potential energy in an electric field described by the Thomas-Fermi potential <f>(r). The second term is the centrifugal energy, in which we put (/+i) 2 instead of /(/+1), since the motion is quasi-classical. Since the total energy of the electron in the atom is negative, it is clear that, if (for given values of Z and /) U t {r) > for all r, there can be no electrons in the atom concerned with the given value of the angular momentum /. If we consider any definite value of /and vary Z, it is found that in fact U^r) > everywhere when Z is sufficiently small. As Z is increased, a value is reached for which the curve of U t (r) touches the axis of abscissae, while for larger Z there is a region where C/ Z (r) < 0. Thus the value of Z at which electrons with the given / appear in the atom is determined by the condition that the curve of £/j(r) touches the axis of abscissae, i.e. by the equations Ufc) = -t+W+Wlr 2 = 0, U t \r) = _^(r)-(/+£) 2 /r* = . Substituting here the expression (70.6) for the potential, we obtain the equations Z*l\{x)lx = (4/3^/3(/+i) 2 /^, Z*l*[x x '{x)- X {x)\lx = -2(4/37r) 2 / 3 (/+i) 2 /* 2 . f In books on chemistry, lutetium is also usually placed with the rare-earth elements. This, how- ever, is incorrect, since the 4/ shell is complete in lutetium; it must therefore be placed in the platinum group, as in Table 4. % As in §70, we use atomic units. 258 The Atom §73 s\s <3 iv. ft f o fel w X Q &3 ^ ^ i> ^ i> i> *S> § CO i. &3 i-> « u vO IM OS >-l S> •* C3 •O \o k* 1 to ^ + X 2 <s 9 g K .O ■*»» s ft o SJ « fel 6 o to o >0 <N 4 ■» Co 3 Til 03 CM CO ^ r° <M «0 eg «o if £ X! o o> <o vO 00 SO + V a o c 5 1 Dividing each side of the second equation by the corresponding side of the first, we find for x the equation X '(x)lx(x) = -1/*, §74 X-ray terms 259 and we then calculate Z from the first of equations (73.2). A numerical calculation gives Z = 0- 155(2/+ 1) 3 . This formula determines the value of Z for which electrons with a given / first appear in the atom; the error is about 10 per cent. Very accurate values are obtained by taking the coefficient as 017 instead of 0155: Z =0-17(2/+ 1) 3 . (73.3) For / = 1, 2, 3 this formula gives respectively, after rounding to the nearest integer, the correct values 5, 21, 58. For / = 4, formula (73.3) gives Z = 124; this means that^ electrons should first appear only in the 124th element. §74. X-ray terms The binding energy of the inner electrons in the atom is so large that, if such an electron makes a transition into an outer unfilled shell (or is removed from the atom), the excited atom (or ion) is mechanically unstable with respect to ionisation, which is accompanied by the reconstruction of the electron envelope and the formation of a stable ion. However, because of the comparatively weak interaction between the electrons in the atom, the prob- ability of such a transition is comparatively small, so that the lifetime t of the excited state is long. Hence the "width" k/r of the level (see §44) is so small that it is reasonable to regard the energies of an atom with an excited inner electron as discrete energy levels of "quasi-stationary" states of the atom. These levels are called X-ray terms.-f The X-ray terms are primarily classified according to the shell from which the electron is removed, or in which, as we say, a "hole" is formed. Where the electron goes has almost no effect on the energy of the atom, and hence is unimportant. The total angular momentum of the set of electrons occupying any shell is zero. When one electron has been removed, the shell acquires some angular momentum j. For the («, /) shell, the angular momentum j can obviously take the values /db£. Thus we obtain levels which might be denoted by ls 1/2 , 2s 1/2 , 2p 1/2 , 2p 3/2 , ..., where the value of j is added as a suffix to the letter giving the position of the "hole". It is usual, however, to employ special symbols as follows : ls 1/2 2s 1/2 2p 1/2 2p3/z 3s 1/2 3p 1/2 3p 3/2 3d 3/2 3d 5/2 ... K Li Lh Liu Mi M u M iu M w My ... The levels with « = 4, 5, 6 are similarly denoted by the letters N, O, P. Levels with the same n (denoted by the same capital letter) lie close together f The name is due to the fact that transitions between these levels cause the emission of X-rays by the atom. 260 The Atom §74 and at a distance from levels with a different n. The reason for this is that, owing to the relative nearness of the inner electrons to the nucleus, they are in the almost unscreened field of the nucleus, and hence their states are "hydrogen-like"; the energy is, to a first approximation, ~Z 2 /2n 2 (in atomic units), i.e. depends only on n. If relativistic effects are taken into account, terms with different j are separated (cf. the discussion in §72 of the fine structure of the hydrogen levels), such as, for example, L\ and Lu from Lm, and Mi and M\\ from Mm and Miy. These pairs of levels are said to be regular (or relativistic) doublets. The separation of terms with different / and the same 7 (for instance L\ and Lu, M\ and Mu) is due to the deviation of the field in which the inner electrons move from the Coulomb field of the nucleus, i.e. to the taking into account of the interaction of the electron with other electrons. These are said to be irregular (or screening) doublets. The main correction term to the "hydrogen- like" energy of the electron results from the potential due to the remaining electrons in the region near the nucleus, and is proportional to Z 4/3 (see (70.8)). However, since this correction does not depend on either n or /, it does not affect the level spacings. The principal correction terms in the level differences are therefore due to the interaction of one electron with those adjoining it. Since the distances between the inner electrons are r ~ 1/Z (the Bohr radius in the field of a charge Z), the energy of this interaction is ~ \jr ~ Z. Taking this correction into account, we can write the energy of an X-ray term, to the same accuracy, as —(Z—8) 2 j2n 2 , where 8 = 8{n,l) is a quantity small compared with Z, and may be regarded as the magnitude of the screening of the nuclear charge. Terms with two and three "holes" may exist in the electron shells together with the X-ray terms with one "hole". Since the spin-orbit interaction is strong for the inner electrons, the holes are subject toji)' coupling. The width of an X-ray term is determined by the total probability of all possible processes by rearrangement of the electron envelope of the atom so as to fill the "hole" in question. In the heavy atoms, transitions of the hole from a given shell to a higher one (i.e. electron transitions in the opposite direction) are the most important, and are accompanied by the emission of X-ray quanta. The probability of these "radiative" transitions, and therefore the corresponding part of the level width, increase very rapidly with the atomic number (as Z 4 ) but decrease towards higher levels for a given Z.f For lighter atoms (and higher levels) an important or even predominant part is played by radiationless transitions, in which the energy liberated when a hole is filled by an electron from above goes to remove another inner electron from the atom (called the Auger effect). As a result of this process the atom is in a state with two holes. The probabilities of these processes and the corresponding contribution to the level width are independent of the atomic number to a first approximation with respect to 1/Z (see the Problem).:): t See Volume 4. X As an example it may be mentioned that the Auger width of the K level is about 1 eV, and reaches values of the order of 10 eV for higher levels. §75 Multipole moments 261 PROBLEM Find the limiting law of dependence of the Auger width of X-ray terms on atomic number when the latter is sufficiently large. Solution. The Auger transition probability is proportional to the square of a matrix element of the form M = ( Uw^VfafcdVidVz, where ipi, <p2 and if/^, tp'a are the initial and final wave functions of the two electrons involved in the transition, and V = e 2 /ri2 is their interaction energy. When Z is sufficiently large, the wave functions of the inner electrons may be regarded as hydrogen-like and the screening of the field of the nucleus by other electrons may be neglected (the wave function of the ionisa- tion electron is also hydrogen-like in the region within the atom which is of importance in the integral M). If we carry out the calculations and all quantities are expressed in Coulomb units (with the constant a = Ze 2 ; see §36), then the only quantity in the integral M which depends on Z is V = \jZriz, so that M ~ \\Z. The transition probability, and therefore the Auger width AE of the level, are proportional to \jZ 2 . On returning to ordinary units (the Coulomb unit of energy being Z 2 me i jh 2 ), we find that AE is independent of Z. §75. Multipole moments In classical theory, the electrical properties of a system of particles are described by its multipole moments of various orders, expressed in terms of the charges and co-ordinates of the particles. In the quantum theory, the definitions of these quantities are the same in form, but they must now be regarded as operators. The first multipole moment is the dipole moment, defined as the vector d= Ser, where the summation is over all the particles, and the suffix which numbers the particles is omitted for brevity. The matrix of this operator, like that of any polar vector (see §30), has non-zero elements only for transitions between states of different parity. The diagonal elements are therefore always zero. In other words, the mean values of the dipole moment of any system of par- ticles (e.g. an atom) in stationary states are zero.f The same is evidently true of all 2*-pole moments with odd /. The com- ponents of such a moment are polynomials of odd degree / in the co-ordinates, which, like the components of a polar vector, change sign on inversion of the co-ordinates. The same parity selection rule therefore applies. The quadrupole moment of a system is defined as the symmetrical tensor Q tk ='Ze(3xiX k -8i k r 2 ), (75.1) t To avoid misunderstanding it should be emphasised that this refers to a closed system of particles or to a system of particles in a centrally symmetric external electric field. For example, if the nuclei are regarded as "fixed", the above statement is valid for the electrons in an atom, but not for those in a molecule. It is also assumed that there is no additional ("accidental") degeneracy of the energy level other than that with respect to directions of the total angular momentum. If this is not so, wave functions of stationary states can be constructed which do not have any definite parity, and the corresponding diagonal elements of the dipole moment need not vanish. 262 The Atom §75 the sum of whose diagonal terms is zero. The determination of the values of these quantities in a particular state of a system (an atom, say) requires an averaging of the operator (75.1) over the corresponding wave function. This averaging is conveniently carried out in two stages (cf. §72). Let Q ile denote the quadrupole moment operator averaged over the electron states with a given value of the total angular momentum / (but not of its com- ponent Mj). The only vector which pertains to the atom as a whole is the "vector" J. The only symmetrical tensor operator with zero trace is therefore of the form IQ Q« = UtJic+Mi-mvc)- (75.2) The constant Q is defined so that Q zz = Q in the state with Mj = J; this quantity is usually called simply the quadrupole moment. The operators J\ must be regarded as the known matrices of the angular momentum with a given value of J but non- diagonal with respect to Mj\ the operator j 2 can, of course, be simply replaced by its eigenvalue /(/+ 1). For / = (so that Mj = also) all the elements of these matrices are zero, i.e. the operators vanish identically. This occurs also when / = \. This is easily seen by direct multiplication of the Pauli matrices (55.6), wh ch are the matrices of the components of any angular momentum equal to£. This is no accident, but is a particular case of the general rule that the tensor of a 2*-pole moment (with even /) is non-zero only for states of the system with total angular momentum / > |/.f PROBLEMS Problem 1. Find the relation between the operators of the quadrupole moment of an atom in states corresponding to various components of the fine structure of a level (i.e. states with different values of J but given values of L and S). Solution. In states with given values of L and S, the operator of the quadrupole moment, a purely "orbital" quantity, depends only on the operator L, and so is given by the same formula (75.2) with J replaced by L and with a different constant Q. The operator (75.2) is obtained by a further averaging over the state with a given value of/: SQj . . Qik = 27(27- 1) W* + -foft-&fl/+ ^ 3Ql 2L(L-1) [til k +t k ti-%L(L+l)8t k ]. (1) t This rule is a consequence of the general properties of symmetry with respect to rotations. The tensor of the 2*-pole moment £)(') is an irreducible tensor of rank / (see The Classical Theory of Fields, §41); its transformation properties correspond to those of a symmetrical spinor of rank 21 (§58). The wave functions t/ij correspond to a spinor of rank 2J. The matrix elements of the operator Z)W are non-zero if, in the integrals which determine them, the integrand (tfu* DW xfij) contains a scalar part. This part is obtained by contraction with respect to all pairs of spinor indices, with each pair belonging to different factors (tfu*, ipj and DM); otherwise the result is zero. It is clear that such a contraction is possible only if the sum of the numbers of indices in tfu and tfu* is not less than the number of indices in £>('), i.e. if 4/ ^ 21. §75 Multipole moments 263 It is required to find the relation between the coefficients Qj and Ql> To do so, we multiply equation (1) on the left by ji and on the right by /*, sum over i and k, and take the eigen- values of the diagonal operators. We have where, by (31.3), 2JX =/(/+l)+L(L+l)-S(S+l). The product ji £* Li j* can be transformed by means of the formulae {LiLic} = ieacitu {Jiti} = ieumLm, as in §29, Problem 2; the result is jit k ij k = (m*-j.L. Similarly MW* = (J 2 ) 2 , AW*/*=J 2 (J 2 -i)- Thus we obtain from (1) the relation 3J.L(2J.L- 1 )-2/(7 + l)L(£ + l) (7+l)(2/+3)L(2i-l) In particular, for S = \ this formula gives Qj=Ql for/ = L+J, (L-l)(2L+3) Qj= Ql- for 7 = L-*. * * L(2L+1) (3) Problem 2. Express the quadrupole moment of an electron (charge —e) with orbital angular momentum / in terms of the mean square of its distance from the centre. Solution. We have to average the expression Q zz = - er\Z cos20- 1) = -er\ln z *-\) over a state with given angular momentum I and component m — I. The mean value of the angle factor is found immediately from the formula derived in §29, Problem 2 (where l z must be replaced by /) ; the result is - 2/ Qi = er* . (4) * 2/+3 The sign of this quantity is opposite to that of the electron charge (— e), as it should be: a particle moving with an angular momentum in the jsr-direction is mainly near the plane z = 0, and hence cos 2 < \. For an electron with a given value of / = / ± £, formulae (3) give & = " 2 (2/-l)/(2/+2). (5) Problem 3. Determine the quadrupole moment of an atom (in the ground state) in which all v electrons in excess of closed shells are in equivalent states with orbital angular momen- tum /. 264 The Atom §75 Solution. Since the total quadrupole moment of completed shells is zero, the quad- ruple moment operator of the atom is given by the sum Qa = (2/-1X2/+3) 2 |> +/ ' ? '-^+ D«4 taken over the v outer electrons (here we have used formula (4)). Let us first suppose that v ^ 21+1, i.e. at most half the places in the shell are occupied. Then, by Hund's rule (§67), the spins of all the v electrons are parallel (so that S = £v). This means that the spin wave function of the atom is symmetrical, and the co-ordinate wave function therefore antisymmetrical, with respect to these electrons. Thus the electrons must all have different values of m, so that the greatest possible value of Ml (and the value of L, which is the same) is I L = (M L ) m&x = SJ m = \v{2l-v+\). The required Ql is the eigenvalue Q zz for Ml = L. We therefore have 6er 2 (2/-l)(2/+3) 6er 2 ^-v (2/-1Y2/+3) ^ L 3V }b whence, on calculating the sum, 2/(2/- 2v+\) _ Ql = — -er\ (2/-l)(2/+3) The final change from Ql to Qj is effected by means of formula (2). The case of an atom whose outer shell is more than half filled is reduced to the previous one by considering "holes" instead of electrons: the result is therefore given by the same formula (6) with the opposite sign (the "hole" charge being +e), v being now taken not as the number of electrons but as the number of unoccupied places in the shell. Problem 4. Determine the quadrupole splitting of the levels in an axially symmetric external electric field.f Solution. Owing to the axial symmetry of the field (taking the axis of symmetry to be the 2-axis) we have d^/dx 2 = d^/dy 2 , where <f> is the field potential, and A<f> = from the electrostatic equation; hence d^/dz 2 = -28m dx 2 . The Hamiltonian of the quadrupole moment in the external field isf #=£ 3 a2 + 1> whence in this case 8=AJx+L 2 )-2AJ z 2 = A{p-ZJ z % "2/(2/- 1) dx*. Replacing the operators by their eigenvalues, we obtain the displacement of the levels: AE = A[J(J+l)-3Mj2]. t A similar problem for an arbitrary field is discussed in §103, Problem 5. t See The Classical Theory of Fields, §42. §76 The Stark effect 265 §76. The Stark effect If an atom is placed in an external electric field, its energy levels are altered ; this phenomenon is known as the Stark effect. In an atom placed in a homogeneous external electric field, we have a system of electrons in an axially symmetric field (the field of the nucleus together with the external field). The total angular momentum of the atom is therefore, strictly speaking, no longer conserved; only the projection Mj of the total angular momentum J on the direction of the field is conserved. The states with different values of Mj have different energies, i.e. the electric field removes the degeneracy with respect to directions of the angular momentum. The removal is, however, incomplete: the states differing only in the sign of Mj are degenerate as before. For an atom in a homogeneous external electric field is symmetrical with respect to reflection in any plane passing through the axis of symmetry (i.e. the axis passing through the nucleus in the direction of the field; we shall take this as the sr-axis). Hence the states obtained from one another by such a reflection must have the same energy. On reflection in a plane passing through some axis, however, the angular momentum about this axis changes sign (the direction of a positive revolution about the axis becomes that of a negative one). We shall suppose that the electric field is so weak that the additional energy due to it is small compared with the distances between neighbouring energy levels of the atom, including the fine-structure intervals. Then, in order to calculate the displacement of the levels in the electric field, we can use the perturbation theory developed in §§38 and 39. Here the perturbation operator is the energy of the system of electrons in the homogeneous field <f, and this is V= -d. «?= -£d 2 , (76.1) where d is the dipole moment of the system. In the zeroth approximation, the energy levels are degenerate (with respect to directions of the total angular momentum); in the present case, however, this degeneracy is unimportant, and in applying perturbation theory we can proceed as if we were dealing with non- degenerate levels. This follows from the fact that, in the matrix of the quantity d z (as in that of the ^-component of any vector), only the elements for transitions without change of Mj are not zero (see §29), and hence states with different values of Mj behave independently when perturbation theory is applied. The displacement of the energy levels is determined, in the first approxi- mation, by the diagonal matrix elements of the perturbation. But all the diagonal matrix elements of the dipole moment vanish identically (§75). Thus the splitting of the levels in an electric field is a second-order effect and is proportional to the square of the field.f It must be calculated according f The hydrogen atom forms an exception; here the Stark effect is linear in the field (see the next section). The atoms of other elements, when in highly excited states (and therefore hydrogen-like; see §68), behave like hydrogen in sufficiently strong fields. 266 The Atom §76 to the general rules of perturbation theory, but the dependence of the dis- placement of the levels on the quantum number Mj can be established from more general considerations. Being quadratic in the field, the displacement AE n of the level E n must be of the form AE n = - £a a (n) <W*, (76.2) where a^") is a symmetrical tensor of rank two; taking the s-axis in the direction of the field, we obtain AE n = -\^n)^. (7 6 .3) The tensor a tt <»> is specific to the (unsplit) level concerned, and depends also on the number Mj. The values of a tt <»> for various values of Mj can be regarded as eigenvalues of the operator a« (B) = KnSvc+pniJiJk+JicJi-iSap). (76.4) This is the most general symmetrical tensor of rank 2 depending on the vector j (cf. §75). From (76.3) and (76.4) we have AE n = -K 2 K+2£„[M./2-i/C/ + 1)]}. (76.5) It may be noted that, on summation over all values of Mj t the second term in the braces vanishes, so that the first term is the displacement of the "centre of gravity" of the split level. Moreover, according to (76.5) a level with J =\ remains unsplit, in accordance with Kramers' theorem (§60). The tensor a«f<») which appears in the above formulae is also the polaris- ability of the atom in the external electric field. According to a general formula, (8Hld\) nn = dE n ld\ (see §11, Problem). Taking the parameters A to be the components of the vector € and putting fi = i? -^.d, we find for the mean value of the dipole moment of the atom dt = a^Hf*. (76.6) If the atom is in a non-uniform external field (which varies only slightly over the dimensions of the atom), there can also exist a splitting effect linear in the field, due to the quadrupole moment of the atom. The operator of the quadrupole interaction between the system and the field has the form which corresponds to the classical expression-)- for the quadrupole energy: ^ =i ^:^' (76 - 7) f See The Classical Theory of Fields, §42. 76§ The Stark effect 267 where <f> is the potential of the electric field (the derivatives being understood to be taken at the position of the atom). We can, in particular, apply this formula to a neutral atom in the field of an electron which is at a distance large compared with atomic dimensions. Then the field of the electron at the position of the atom satisfies the con- dition of quasi-uniformity, and in the first approximation of perturbation theory we find that the energy of interaction between the electron and the atom is proportional to 1/r 3 (since j> ~ 1/r). This result applies, however, only to states with given values of the com- ponent Mj of the total angular momentum. On averaging over all directions of the angular momentum (i.e. over all values of Mj\ the interaction pro- portional to 1/r 3 vanishes, since Qu = 0. Furthermore, this interaction does not exist if the angular momentum J of the atom is or |, since then Qm = ; see §75. The next term in powers of 1/r, which is never zero, is the interaction which appears in the second-order perturbation theory with respect to the dipole operator (76.1). Since the field of the electron & ~ 1/r 2 , the energy U of this interaction is proportional to 1/r 4 . If the atom is in the normal state, this energy (like any second-order correction to the energy of the ground state ; see §38) is negative, i.e. there is a force of attraction between the atom and the electron. This attraction is the reason why some atoms are able to form negative ions by the attachment of an electron.! This property is not shared by all atoms, however, the reason being that, in a field which decreases at large distances as 1/r 4 (or 1/r 3 ), the number of levels (corresponding to a finite motion of the electron) is always finite, and in some cases there may be no such levels. PROBLEMS Problem 1. Determine the Stark splitting of the different components of a multiplet level as a function of /. Solution. The problem is conveniently solved by changing the order in which the per- turbations are applied ; we first consider the Stark splitting of the level in the absence of fine structure, and then bring in the spin-orbit interaction. Since the spin of the atom does not interact with the external electric field, the Stark splitting of a level with orbital angular momentum L is given by a formula of the same form (76.2), with a tensor a.ac which is ex- pressed in terms of the operator L in the same way as oeifc in (76.4) is expressed in terms of J: «4k = a8ik+b(LiLjc+LicLi—%8ikL 2 ), the suffixes n being everywhere omitted. When the spin-orbit interaction is included, the states of the atom must be described by the total angular momentum J. The averaging of the operator oloc over states with a given value of the angular momentum J (but not of its component Mj) is formally identical with the averaging carried out in §75, Problem 1. We hus return to formulae (76.4), (76.5), with constants a,0 which are given in terms of the f For example, the halogen atoms attach an electron with a binding energy of the order of 2 to 4 eV, and the hydrogen atom does so with a binding energy 0*7 eV. 268 The Atom §76 constants a, b by 3J.L[2J.L-l]-2/(7+l)L(L+l) a = a, p = b — . 7(/+l)(2/-l)(2/+3) This determines the splitting as a function of J (but not of L and S, of course; these are characteristics of the unsplit term on which the constants a and b also depend). Problem 2. Determine the splitting of a doublet level (spin S = £) in an arbitrary (not weak) electric field. Solution. If the splitting is not small in comparison with the interval between the components of the doublet, the perturbation from the electric field and the spin-orbit inter- action must be taken into account simultaneously, i.e. the perturbation operator is the sum P = AS.L-$<?2{a+2b[L z 2-±L(L+l)]} (cf. (72.4) and Problem 1). Omitting the constant terms which do not affect the splitting, we can write this operator in the form V = \A[S + L-+ S-L + +2£ z l z -b^L z z] (see (29.11)). For each given value of Mj the eigenvalues of this operator are determined by the roots of the secular equation formed from its matrix elements with respect to the states : (2)M L = Mj+l M s = -%. From formulae (27.12) we find Vn = \A(Mj-\)-b£\Mj-\)\ V22 = -lA(Mj+i)-b£2(Mj+i)2, V 12 = iAV[(L+Mj+$)(L-Mj+$]. Thus (see §39, Problem 1) the level displacement is AE= -bg2M J 2±V[lA2(L+i)2+b<?2(b£ , 2+A)Mj2], (1) where all terms which are the same for all components of the split doublet are omitted. This formula (with both signs of the root) applies to all levels with \Mj\ < L— J. For Mj = L + \ {or Mj = — (L + i)) there is no state 2 (or 1); for this level the displacement is given simply by the matrix element Vn (or F 22 ), i.e., with the same choice of the additive constant as in (1), AE = {\A+b£*){L+\)-b£Z{L+®z. (2) This is the same as the result obtained from formula (1) with only one sign of the root. Problem 3. Determine the quadrupole splitting of levels in an axially symmetric electric field. Solution. In a field symmetrical about the s-axis we have 2 «£/&e 2 = 8 2 <f>/dy 2 = a, 8^<f>ldz 2 = —2a, the remaining second derivatives being zero. The quadrupole energy operator (76.7) is a - - - Qo<* „ -iQxx+Qyy-2Q zz )= (J 2 -3/ 2 2). 6 2/(2/- 1) Replacing the operators by their eigenvalues, we obtain for the displacement of the levels AE = a [J(J+ 1) - 3Mj2]. §77 The Stark effect in hydrogen 269 §77. The Stark effect in hydrogen The levels of the hydrogen atom, unlike those of other atoms, undergo a splitting proportional to the field (the linear Stark effect) in a uniform electric field. This is due to the occurrence of an accidental degeneracy in the hydrogen terms, whereby states with different / (for a given principal quantum number n) have the same energy. The matrix elements of the dipole moment for transitions between these states are not zero, and hence the secular equation gives a non-zero displacement of the levels, even in the first approxi- mation. For purposes of calculation-}- it is convenient to choose the unperturbed wave functions so that the perturbation matrix is diagonal with respect to each group of mutually degenerate states. It is found that this is achieved by quantising the hydrogen atom in parabolic co-ordinates. The wave func- tions ift nintm of the stationary states of the hydrogen atom in parabolic co- ordinates are given by formulae (37.15) and (37.16). The perturbation operator]: (the energy of an electron in the field &) is Sz = \ &($—-n)> the field is directed along the positive s-axis, and the force on the electron along the negative sr-axis. We are interested in the matrix elements for transitions n^m -> n^n^m', for which the energy (i.e. the prin- cipal quantum number n) is unaltered. It is easy to see that, of these, only the diagonal matrix elements CO0O27T J ^..J 1 '* dV = K jjj (e~V 2 )\Kn,m\ 2 d^drj = K fjfn, m 2 (Pl)fn 2 m 2 (P2)(Pl 2 -P2 2 ) <WP2 (77-1) are non-zero (we have made the substitution £ = np v t] = »p 2 ). The matrix concerned is evidently diagonal with respect to the number m, while its diagonality with respect to the numbers n 1 , n 2 follows from the orthogonality of the functions f niTn for different n x and the same n (see below). The integra- tions over p x and p 2 in (77.1) are separable; the integrals obtained are calcu- lated in §f of the Mathematical Appendices (integral (f.6)). After a simple calculation, we find for the corrections to the energy levels in the first approxi- mation || £(« = | £n{n x -n 2 \ (77.2) or, in absolute units, f In the following calculations we do not take account of the fine structure of the hydrogen levels. Hence the field must be, though not strong (for perturbation theory to be applicable), yet such that the Stark splitting is large in comparison with the fine structure. % In this section we use atomic units. || This result was derived by K. Schwarzschild and P. Epstein (1916), using the old quantum theory, and by W. Pauli and E. Schrodinger (1926) using quantum mechanics. 270 The Atom §77 The two extreme components of the split level correspond to Wj = n— 1, «2 = and % = 0, w 2 = n— 1. The distance between these two extreme levels is, by (77.2), Z£n{n-\\ i.e. the total splitting of the level by the Stark effect is approximately pro- portional to n 2 . It is natural that the splitting should increase with the prin- cipal quantum number: the further the electrons are from the nucleus, the greater the dipole moment of the atom. The presence of the linear effect means that, in the unperturbed state, the atom has a dipole moment whose mean value is d~ e = -fw^-^). (77.3) This is in accordance with the fact that, in a state determined by parabolic quantum numbers, the distribution of the charges in the atom is not sym- metrical about the plane z = (see §37). Thus, for % > n^, the electron is predominantly on the side of positive z, and hence the atom has a dipole moment opposite to the external field (the charge on the electron being negative). In the previous section we have shown that a uniform electric field cannot entirely remove degeneracy: there always remains a twofold degeneracy of states differing in the sign of the projection of the angular momentum on the direction of the field (in this case, states whose projected angular momenta are + m). However, we see from formula (77.2) that even this removal of the degeneracy does not occur in the linear Stark effect in hydrogen: the displacement of the levels (for given n and n± - n 2 ) is independent of m and n 2 . A further removal of the degeneracy occurs in the second approximation; the calculation of this effect is the more interesting in that the linear Stark effect is altogether absent in states with n\ = w 2 - To calculate the quadratic effect, it is not convenient to use ordinary perturbation theory, since it would be necessary to deal with infinite sums of complicated form. Instead we use the following slightly modified method. Schrodinger's equation for the hydrogen atom in a uniform electric field is of the form (£A+£+i/r-«f*)«A = o. Like the equation with «f = 0, it allows separation of the variables in parabolic co-ordinates. The same substitution (37.7) as was used in §37 gives the two equations d / d/„ \ / m 2 \ d^Vd^J + (^ _i ~ +i<f V /2 = ~^ 2y (77.4) &+& = 1, §77 The Stark effect in hydrogen 271 which differ from (37.8) by the presence of the terms in S. We shall regard the energy E in these equations as a parameter which has a definite value, and the quantities jS 1} j3 2 as eigenvalues of corresponding operators; it is easy to see that these operators are self-conjugate. These quantities are deter- mined, by solving the equations, as functions of E and S, and then the condi- tion j8i+j82 = 1 gives the required relation between E and S, i.e. the energy as a function of the external field. For an approximate solution of equations (77.4), we regard the terms con- taining the field ^ as a small perturbation. In the zeroth approximation ( g = 0), the equations have the familiar solutions A = Ve/n^e), (77.5) /■ = V e fn t mM, where the functions f nitn are the same as in (37.16), and instead of the energy we have introduced the parameter € = V(-2E). ( 776 > The corresponding values of &, jS 2 (from the equations (37.12), in which n must be replaced by 1/e) are &(« = K+iM+Jh j8 2 «» = ( W2 +i|m|+£)e. (77.7) The functions j x with different % for a given e are orthogonal, as are the eigen- functions of any self-conjugate operator; we have already used this fact above in discussing the linear effect. In (77.5) these functions are normalised by the conditions o o The corrections to & and jS 2 in the first approximation are determined by the diagonal matrix elements of the perturbation: oo oo Calculation gives &W = i^(6n 1 2 +6n 1 |m|+m 2 +6n 1 +3|m|+2)/€ 2 . The expression for j8 2 (1) is obtained by replacing n x by n^ and changing the sign. 272 The Atom §77 In the second approximation we have, by the general formulae of perturba- tion theory, <f 2 ^ Kaw The integrals appearing in the matrix elements (£ 2 )„ „ . are calculated in S f of the Mathematical Appendices. The only non-zero elements are (^ 2 )« 1 ,« 1 -i = (P ) ni -i,n = -2(2« 1 +M)V[h 1 (« 1 +M)]/ 6 2, (?)n it n r 2 = {¥\-2,n x = VM^- 1)K+ |m|)K+|m|- 1)]{ € *. The differences occurring in the denominators are As a result of the calculations we have ft(2> = - < r 2 (|m|+2 % +l)[4m 2 +17(2|m| % +2« 1 H|M+2n 1 )+18]/1665; the expression for j8 2 < 2 > is obtained by replacing « x by n^. Combining the expressions obtained and substituting in the relation ft+/? 2 = 1, we have the equation en- < r 2 «[17n 2 +51K~« 2 )2-9m 2 +19]/16 € 5+3 ( f w ( Wi _ M2 )/ € 2 = x Solving by successive approximations, we have in the second approximation for the energy E = — \e % the expression 1 «f 2 E = ~2^ 2 " +l ^" (ni ~ M2) ~l6" w4[17w2 ~ 3(Wl ~ W2)2 ~ 9m2+19] ' (77 * 8) The second term is the already familiar linear Stark effect, and the third is the required quadratic effect (G. Wentzel, I. Waller and P. Epstein 1926). We notice that this quantity is always negative, i.e. the terms are always displaced downwards by the quadratic effect. The mean value of the dipole moment is obtained by differentiating (77.8) with respect to the field; in the states with n\ = n^ it is d z =£M 4 (17» 2 -9m 2 +19)«f. (77.9) Thus the polarisability of the hydrogen atom in the normal state (»=1, m=0) is 9/2 (in absolute units 9(^ a /we 2 ) 3 /2). The absolute value of the energy of the hydrogen terms falls rapidly as the principal quantum number n increases, while the Stark splitting is increased. Hence it is of interest to examine the Stark effect for highly excited levels in §77 The Stark effect in hydrogen 273 fields so strong that the splitting they cause is comparable with the energy of the level itself, and perturbation theory is inapplicable.! This can be done by using the fact that states with large values of n are quasi-classical. By the substitution A=xJVe, h=xM (77-10) the equations (77.4) are brought into the form / ft w 2 -l a \ (77.11) drf Each of these equations, however, is the same in form as the one-dimensional Schrodinger's equation, the part of the total energy of the particle being taken by \E, and that of the potential energy by the functions ft m 2 -l ft w 2 -l (77.12) 2tj Sr, 2 respectively. Figs. 25 and 26 respectively show the approximate form of these functions (for m > 1). By Bohr and Sommerfeld's quantisation rule (48.2) we write i, j y/{2&E-Ujm d£ = K+iK ix (77.13) VWhE-UtivWdrt =(« 2 +|K where %, w 2 are integers. J These equations determine implicitly the depen- dence of the parameters & and jS 2 on E. Together with the equation ft + 2 = 1 , they therefore give the energies of the levels when displaced by the electric field. The integrals in equations (77.13) can be reduced to elliptic integrals; these equations can be solved only numerically. The applicability of perturbation theory to high levels requires the perturbation to be small only in comparison with the energy of the level itself (the binding energy of the electron), and not with the intervals between the levels. For in the quasi-classical case (which corresponds to highly excited states) the perturbation can be regarded as small if the force due to it is small in comparison with those acting on the particle in the unperturbed system; and this condition is equivalent to the one given above. % A detailed investigation shows that a more exact result is obtained by writing m z instead of m'— 1 in the expressions for U x and U 2 . The integers n lt Wj are then equal to the parabolic quantum numbers. 274 The Atom §77 Fig. 25 The Stark effect in strong fields is complicated by another phenomenon, the ionisation of the atom by the electric field. The potential energy £z of an electron in the external field takes arbitrarily large negative values as z -» — co. Added to the potential energy of the electron within the atom, it has the effect that the region of possible motion for the electron (whose total energy E is negative) includes, besides the region inside the atom, the region of large distances from the nucleus in the direction of the anode. These two regions are separated by a potential barrier, whose width dimin- ishes as the field increases. We know, however, that in quantum mechanics there is always a certain non-zero probability that a particle will penetrate a potential barrier. In the case we are considering, the emergence of the elec- tron from the region within the atom, through the barrier, is simply the ionisation of the atom. In weak fields the probability of this ionisation is vanishingly small. It increases exponentially with the field, however, and becomes considerable in fairly strong fields.-f Fig. 26 t This phenomenon may serve to illustrate how a small perturbation may alter the nature of the energy spectrum. Even a weak field $ is sufficient to create a potential barrier and produce a region, far from the nucleus, which is in principle accessible to the electron. As a result, the motion of the electron becomes, strictly speaking, infinite, and hence the energy spectrum becomes continuous instead of discrete. Nevertheless, the formal solution obtained by the methods of perturbation theory has a physical significance: it gives the energy levels of states which are not quite but "almost" station- ary. An atom that is in such a state at some initial instant remains in it for a long period of time. However, the series given by perturbation theory for the Stark splitting of the levels cannot be convergent in the strict sense, but is merely an asymptotic series : after a certain point in the series (which becomes later as the perturbation is reduced in magnitude) the terms increase, not decrease. §77 The Stark effect in hydrogen 275 PROBLEM Determine the probability (per unit time) of the ionisation of a hydrogen atom (in the ground state) in a strong electric field. Solution. In parabolic co-ordinates there is a potential barrier "along the 17 co-ordinate" (Fig. 26); the "extraction" of the electron from the atom in the direction z -> — 00 corresponds to its passage into the region of large 17. To determine the ionisation probability, it is necessary to investigate the form of the wave function for large 17 (and small £ ; we shall see below that small values of £ are the important ones in the integral which determines the total probability current for the emerging electron). The wave function of the electron in the normal state (in the absence of the field) is = «-tf+W2/yV. (1) When the field is present, the dependence of iff on £ in the region in which we are interested can be regarded as being the same as in (1), while to determine its dependence on 77 we have the equation d\ r 11 1 v + [- i+ ^ + v +K, ] x = 0, (2) where \ =■ y/-r\i\> (equation (77.11) with E = — £, m = 0, /3 2 = £). Let 17 be some value of 17 ("within" the barrier) such that 1 <^ 7? <C l/<^. For 17 <; 179, the wave function is quasi- classical. Since, on the other hand, equation (2) has the form of the one-dimensional Schrodinger's equation, we can use formulae (50.4). Using the exact expression (1) for the wave function at the point 17 = 170, we obtain in the region outside the barrier the expres- sion 1 x l\Po\ } X = — -€ r -* ( * f "«V 1 7o / ex p[* \pdr)— f«r], where We shall be interested only in the square |x| 2 - Hence the imaginary part of the exponent is unimportant. Denoting by -q x the root of the equation p(rj) = 0, we have | x |2 = lV^exp[-2f|/> |d,-, ]. 71 P i In the coefficient of the exponential we put in the exponent we must keep also the next term of the expansion : Vi Vi \x\ 2 = er€ exp[~- f y/ll-g-q) d v + f « "|, where ^ ^ \\S. Effecting the integration and neglecting 17 <^ compared with 1 wherever possible, we obtain |x| 2 = e- 2 /^ . (3) TtS \Z(<a7)—l) 276 The Atom §77 The total probability current through a plane perpendicular to the ar-axis (i.e. the required ionisation probability w) is 00 w = \*fi\H z 2TTr dr. o For large rj (and small £) we can put Substituting also for the velocity of the electron v. s V[2(-l+*^)] = VW-1), we have 00 W=j\x\ 2 7rV(<?ri-l)d$ o 4 °? ^ o that is, finally, or, in ordinary units, w = (4m3<?9/«f/z7)exp {-2m^\ZSh% CHAPTER XI THE DIATOMIC MOLECULE §78. Electron terms in the diatomic molecule In the theory of molecules an important part is played by the fact that the masses of atomic nuclei are very large compared with those of the electrons. Because of this difference in mass, the rates of motion of the nuclei in the molecule are small in comparison with the velocities of the electrons. This makes it possible to regard the motion of the electrons as being about fixed nuclei placed at given distances from one another. On determining the energy levels U n for such a system, we find what are called the electron terms for the molecule. Unlike those for atoms, where the energy levels were certain numbers, the electron terms here are not numbers but functions of parameters, the distances between the nuclei in the molecule. The energy U n includes also the electrostatic energy of the mutual interaction of the nu- clei, so that U n is essentially the total energy of the molecule for a given arrangement of the fixed nuclei. We shall begin the study of molecules by taking the simplest type, the diatomic molecules, which permit the most complete theoretical investigation. One of the chief principles in the classification of the atomic terms was the classification according to the values of the total orbital angular momen- tum L. In molecules, however, there is no law of conservation of the total orbital angular momentum of the electrons, since the electric field of several nuclei is not centrally symmetric. In diatomic molecules, however, the field has axial symmetry about an axis passing through the two nuclei. Hence the projection of the orbital angular momentum on this axis is here conserved, and we can classify the electron terms of the molecules according to the values of this projection. The absolute value of the projected orbital angular momentum along the axis of the molecule is customarily denoted by the letter A; it takes the values 0, 1, 2, ... . The terms with different values of A are denoted by the capital Greek letters corresponding to the Latin letters for the atomic terms with various L. Thus, for A = 0, 1, 2 we speak of S, II and A terms respectively; higher values of A usually need not be considered. Next, each electron state of the molecule is characterised by the total spin S of all the electrons in the molecule. If S is not zero, there is degeneracy of degree 2S+1 with respect to the directions of the total spin. f The number 2S+1 is, as in atoms, called the multiplicity of the term, and is written as an index before the letter for the term; thus 3 II denotes a term with A = 1 S=l. f We here neglect the fine structure due to relativistic interactions (see §§83 and 84 below). 277 278 The Diatomic Molecule §78 Besides rotations through any angle about the axis, the symmetry of the molecule allows also a reflection in any plane passing through the axis. If we effect such a reflection, the energy of the molecule is obviously unchanged. The state obtained from the reflection is, however, not completely identical with the initial state. For, on reflection in a plane passing through the axis of the molecule, the sign of the angular momentum about this axis is changed. Thus we conclude that all electron terms with non-zero values of A are doubly degenerate: to each value of the energy, there correspond two states which differ in the direction of the projection of the orbital angular momentum on the axis of the molecule. In the case where A = the state of the molecule is not changed at all on reflection, so that the S terms are not degenerate. The wave function of a S term can only be multiplied by a constant as a result of the reflection. Since a double reflection in the same plane is an iden- tity transformation, this constant is ±1. Thus we must distinguish S terms whose wave functions are unaltered on reflection and those whose wave functions change sign. The former are denoted by S+, and the latter by S~. If the molecule consists of two similar atoms, a new symmetry appears, and with it an additional characteristic of the electron terms. A diatomic molecule with identical nuclei has a centre of symmetry at the point bisecting the line joining the nuclei.f (We shall take this point as the origin.) Hence the Hamiltonian is invariant with respect to a simultaneous change of sign of the co-ordinates of all the electrons in the molecule (the co-ordinates of the nuclei remaining unchanged). Since the operator of this transformation J also commutes with the orbital angular momentum operator, we have the possibility of classifying terms with a given value of A according to their parity: the wave functions of even (g) states are unchanged when the co- ordinates of the electrons change sign, while those of odd (u) states change sign. The suffixes u, g indicating the parity are customarily written with the letter for the term: Ii u , Rg, and so on. Finally, we shall mention an empirical rule, according to which the normal electron state in the great majority of chemically stable diatomic molecules is completely symmetrical: the electron wave function is invariant with respect to all symmetry transformations in the molecule. As we shall show in §81, the total spin S is zero too, in the great majority of cases, in the normal state. In other words, the ground term of the molecule is X S+, and it is 1 E+ i? if the molecule consists of two similar atoms. || PROBLEM Effect the separation of variables in Schrodinger's equation for the electron terms of the ion H2 + , using elliptic co-ordinates. f It has also a plane of symmetry perpendicularly bisecting the axis of the molecule. This element of symmetry need not be considered separately, however, since the existence of such a plane follows automatically from the existence of a centre of symmetry and of an axis of symmetry. % Not to be confused with that of inversion of the co-ordinates of all the particles in the molecule || Exceptions to these rules are formed by the molecules 2 (whose normal term is 3 S~ g ) and NO (normal term 2 II). §79 The intersection of electron terms 279 Solution. Schrodinger's equation for an electron in the field of two protons at rest is (using atomic units) ^+2(e+— +— y =o- \ r\ r 2 / n r Zi The elliptic co-ordinates £, 17 are defined by £ = (n + r 2 )IR, v = (rz-r^R; 1<£<oo, -1<i?<1, and the third co-ordinate <£ is the angle of rotation about an axis passing through the two nuclei at a distance R apart.f The Laplacian operator in these co-ordinates is 4 rd 8 8 8n A = — (^ 2 — 1) 1 (1 — t? 2 ) — + 4 r8 8 8 8~[ 1 8* \ = R\{ Putting |J2(|2_1)(1_^2) 0^2 = X(g)Y(riy**, we obtain for X and Y the equations dr dX~] / A2 \ dr dY~] ( A2 \ i 1 -' !2) ^] + (-^ v ^-w) y = ' where A is the separation parameter. Each electron term is described by three quantum numbers : A, and two numbers ng , tin which determine the number of zeros of the functions X(£) and Y(-q). Since all these numbers are related to functions of different variables, there is in general nothing to prevent the inter- section of terms E(R) having different values of any two quantum numbers, including the pair ng , «tj with the same A, even though such terms have the same symmetry (see the footnote to §79). §79. The intersection of electron terms The electron terms in a diatomic molecule are functions of a single para- meter, the distance r between the nuclei. They can be represented graphically by plotting the energy as a function of r. It is of considerable interest to examine the intersection of the curves representing the different terms. Let U x {r), U 2 (r) be two different electron terms. If they intersect at some point, then the functions U 1 and U 2 will have neighbouring values near this point. To decide whether such an intersection can occur, it is convenient to put the problem as follows. Let us consider a point r where the functions U x (r), U%(r) have very close but not equal values (which we denote by E v E 2 ), and examine whether or not we can make U x and U 2 equal by displacing the point a short distance 8r. The energies E x and E 2 are eigenvalues of the Hamiltonian ff of the system of electrons in the field of the nuclei, which are at a distance r from each other. If we add to the distance r an increment f See Mechanics, §48. 280 The Diatomic Molecule §79 8r, the Hamiltonian becomes fi Q +1/, where V = 8r . dffjdr is a small cor- rection ; the values of the functions U lt U 2 at the point r + 8r can be regarded as eigenvalues of the new Hamiltonian. This point of view enables us to determine the values of the terms U x (r), U 2 (r) at the point r +8r by means of perturbation theory, V being regarded as a perturbation to the operator The ordinary method of perturbation theory is here inapplicable, however, since the eigenvalues E v E 2 of the energy in the unperturbed problem are very close to each other, and their difference is in general small compared with the magnitude of the perturbation; the condition (38.8) is not fulfilled. Since, in the limit as the difference E 2 —E x tends to zero, we have the case of degener- ate eigenvalues, it is natural to attempt to apply to the case of close eigenvalues a method similar to that developed in §39. Let tf/ v «/r 2 be the eigenfunctions of the unperturbed operator 3 which correspond to the energies E x , E 2 . As an initial zero-order approximation we take, instead of fa and fa themselves, linear combinations of them of the form <A = c x fa+c 2 fa. C 79 - 1 ) Substituting this expression in the perturbed equation &+?)</,= Efa (79.2) we obtain c x {E x +V-E)fa+c z {E 2 +t-E)fa = 0. Multiplying this equation on the left by fa* and 02* in turn, and integrating, we have two algebraic equations : Ci(E 1 + V 11 -E)+c 2 V 12 = 0, CiV n +c t (E t + V 2t -E) = 0, where V ik = j" «^ # ty*fj k dq. Since the operator V is Hermitian, the quantities V n and F 22 are real, while V x% = V 2X *. The compatibility condition for these equations is E 1 +V n -E V u ' V 21 E 2 +V 22 -E whence we obtain after some calculation = 0, E = K£i+£ 2 + Vu+ V tt )±V\WB 1 -E*+ V n - V 22 ?+\ V 12 \*]. (79.4) This formula gives the required eigenvalues of the energy in the first approxi- mation. If the energy values of the two terms become equal at the point r + 8r (i.e. the terms intersect), this means that the two values of E given by formula §79 The intersection of electron terms 281 (79.4) are the same. For this to happen, the expression under the radical in (79.4) must vanish. Since it is the sum of two squares, we obtain, as the condition for there to be points of intersection of the terms, the equations E 1 -E 2 +V 11 -V 22 =0 1 F 12 =0. (79.5) However, we have at our disposal only one arbitrary parameter giving the perturbation V, namely the magnitude Sr of the displacement. Hence the two equations (79.5) cannot in general be simultaneously satisfied (we sup- pose that the functions tf/ lt «/r 2 are chosen to be real, so that V 12 also is real). It may happen, however, that the matrix element V 12 vanishes identically ; there then remains only one equation (79.5), which can be satisfied by a suit- able choice of 8r. This happens in all cases where the two terms considered are of different symmetry. By symmetry we here understand all possible forms of symmetry: with respect to rotations about an axis, reflections in planes, inversion, and also with respect to interchanges of electrons. In the diatomic molecule this means that we may be dealing with terms of different A, differ- ent parity or multiplicity, or (for S terms) £+ and 2~ terms. To prove this statement it is essential that the operator V (like the Hamil- tonian itself) commutes with all the symmetry operators for the molecule: the operator of the angular momentum about an axis, the reflection and in- version operators, and the operators of interchanges of electrons. It has been shown in §§29 and 30 that, for a scalar quantity whose operator commutes with the angular momentum and inversion operators, only the matrix elements for transitions between states of the same angular momentum and parity are non-zero. This proof remains valid, in essentially the same form, for the general case of an arbitrary symmetry operator. We shall not pause to repeat it here, especially since in §97 we shall give another general proof, based on group theory. Thus we reach the result that, in a diatomic molecule, only terms of differ- ent symmetry can intersect, while the intersection of terms of like symmetry is impossible (E. Wigner and J. von Neumann 1929). If, as a result of some approximate calculation, we obtain two intersecting terms of the same symmetry, they are found to move apart on calculating the next approxi- mation, as shown by the continuous lines in Fig. 27.f We emphasise that this result not only is true for the diatomic molecule, but is a general theorem of quantum mechanics ; it holds for any case where the Hamiltonian contains some parameter and its eigenvalues are consequently functions of that parameter. In a polyatomic molecule, the electron terms are functions of not one but several parameters, the distances between the various nuclei. Let s be the number of independent distances between the nuclei ; in a molecule of iV( > 2) atoms, this number is s = 3N— 6 for an arbitrary arrangement of the nuclei. t There is a curious exception to this rule in the case where the problem of determining the electron terms allows a complete separation of the variables (see §78, Problem). 282 The Diatomic Molecule m U(r) Fig. 27 Each term U n {r ly ... , r s ) is, from the geometrical point of view, a surface in a space of s + 1 dimensions, and we can speak of the intersections of these surfaces in manifolds of varying numbers of dimensions, from (intersection in a point) to s— 1. The derivation given above is wholly valid, except that the perturbation V is here determined not by one but by s parameters, the displacements 8r x , ... , 8r s . Even with two parameters, the two equations (79.5) can in general be satisfied. Thus we conclude that, in polyatomic molecules, any two terms may intersect. If the terms are of like symmetry, the intersection is given by the two conditions (79.5), from which it follows that the number of dimensions of the manifold in which the intersection occurs is s—2. If the terms are of different symmetry, on the other hand, there remains only one condition, and the intersection takes place in a mani- fold of s—1 dimensions. Thus for s = 2 the terms are represented by surfaces in a three-dimensional system of co-ordinates. The intersection of these surfaces occurs in lines (s — 1 = 1) when the symmetry of the terms is different, and in points (s — 2 = 0) when it is the same. It is easy to ascertain the form of the surfaces near the point of intersection in the latter case. The value of the energy near the points of intersection of the terms is given by formula (79.4). In this expression the matrix elements V u , V^, V 12 are linear functions of the dis- placements 8r lt 8r 2 , and hence are linear functions of the distances r lt r 2 themselves. Such an equation determines an elliptic cone, as we know from analytical geometry. Thus, near the points of intersection, the terms are represented by the surface of an arbitrarily situated double elliptic cone (Fig. 28). §80. The relation between molecular and atomic terms As we increase the distance between the nuclei in a diatomic molecule, we have in the limit two isolated atoms (or ions). The question thus arises of the correspondence between the electron terms of the molecule and the states of the atoms obtained by moving them apart. This relation is not one-to-one ; if we bring together two atoms in given states, we may obtain a molecule in various electron states. §80 The relation between molecular and atomic terms u 283 Fig. 28 Let us first suppose that the molecule consists of two different atoms. Let the isolated atoms be in states with orbital angular momenta L^, L 2 and spins S lt S 2 , and let Z^ ^ L 2 . The projections of the angular momenta on the line joining the nuclei take the values M 1 = —L v —L x +1, ... , L^ and M 2 = — L 2 , — L 2 +l, ... , L 2 . The absolute value of the sum Mj+ik/a deter- mines the angular momentum A obtained on bringing the atoms together. On combining all possible values of M x and M 2 , we find the following values for the numbers of times that we obtain the various values of A = jiW^+Mjl : A = L x +L 2 L x +L 2 -l twice four times L x —L 2 Lx-£ 2 -l 2(2L 2 +1) times 2(2L 2 +1) times 2(2L 2 +1) times 2L 2 +1 times. Remembering that all terms with A ^ are doubly degenerate, while those with A = are not degenerate, we find that there will be 1 term with A = L x -\-L 2y 2 terms with A = L 1 -\-L 2 — 1, 2L 2 +1 terms with A = L x — L 2 2L 2 +1 terms with A = L 1 —L 2 —l, 2L 2 +1 terms with A = 0; (80.1) 284 The Diatomic Molecule §80 in all, (2L 2 +1)(L 1 + 1) terms with values of A from to L x -\-L 2 . The spins S lt S 2 of the two atoms combine to form the total spin of the molecule in accordance with the general rule for the addition of angular momenta, giving the following possible values of S: S = S 1 +S 2 , S^Sz-l, ..., | Si-Sal . (80.2) On combining each of these values with each value of A in (80.1), we obtain the complete list of all possible terms in the molecule formed. For 2 terms there is also the question of sign. This is easily resolved by noticing that the wave functions of the molecule can be written, as r -» oo, in the form of products (or sums of products) of the wave functions of the two atoms. An angular momentum A = can be obtained either by adding two non-zero angular momenta of the atoms such that M x = — M 2 , or from M 1 = M 2 = 0. We denote the wave functions of the first and second atoms b y «A (1) m 1 > ^ (2 W For M = \M X \ = \M 2 \ # 0, we form the symmetrised and antisymmetrised products A reflection in a vertical plane (i.e. one passing through the axis of the mole- cule) changes the sign of the projection of the angular momentum on the axis, so that ifj a) M , «A (2) m are changed into ^ a) - M > <A (2) -m respectively, and vice versa. The function \Js + is thereby unchanged, while ifs~ changes sign; the former therefore corresponds to a 2+ term and the latter to a 2~ term. Thus, for each value of M, we obtain one 2+ and one 2~ term. Since M can take L% different values (M = 1, ... , L 2 ), we have in all L 2 2+ terms and L 2 2~ terms. If, on the other hand, M 1 = M 2 = 0, the wave function of the molecule is of the form j/f = ^f (1) ^ (2) - In order to ascertain the behaviour of the func- tion ?/r (1) on reflection in a vertical plane, we take a co-ordinate system with its origin at the centre of the first atom, and the #-axis along the axis of the molecule, and we notice that a reflection in the vertical xs-plane is equivalent to an inversion with respect to the origin, followed by a rotation through 180° about the ^-axis. On inversion, the function (1) o is multiplied by I lt where I x = Jb 1 is the parity of the given state of the first atom. Next, the result of applying to the wave function the operation of an infinitely small rotation (and therefore that of any finite rotation) is entirely determined by the total orbital angular momentum of the atom. Hence it is sufficient to consider the particular case of an atom having one electron, with orbital angular momentum / (and a .sr-component of the angular momentum m = 0) ; on putting L in place of / in the result, we obtain the required solution for any atom. The angular part of the wave function of an electron with m = is, apart from a constant coefficient, P/(cos 6) (see (28.7)). A rotation through 180° about the j-axis is the transformation x -> — x, y -> j, z -> —z or, in §80 The relation between molecular and atomic terms 285 spherical polar co-ordinates, r -+r, 6 -+n— 0, <£->7r — (f>. Then cos 6 -> — cos 6, and the function Pj(cos 6) is multiplied by (— 1)*. Thus we conclude that, as a result of reflection in a vertical plane, the func- tion «^ (1) is multiplied by ( — V) Ll I x . Similarly, i/j (2) is multiplied by ( — 1) L 2 / 2 , so that the wave function «/r = «A (1) «A (2) is multiplied by ( — l) Ll+L2 7 1 7 2 . The term is 2+ or 2~ according as this factor is +1 or — 1. Summarising the results obtained, we find that, of the total number 2L 2 +1 of S terms (each of the appropriate multiplicity), L 2 +l terms are 2+ andL 2 are 2", if ( - 1) L > +Lt I 1 I 2 = + 1 , and vice versa if ( - l) Ll +L *IJ 2 = - 1 . Let us now turn to a molecule consisting of similar atoms. The rules for the addition of the spins and orbital angular momenta of the atoms to form the total S and A for the molecule remain the same here as for a molecule composed of different atoms. The difference is that the terms may be even or odd. Here we must distinguish two cases, according as the combined atoms are in the same or different states. If the atoms are in different states, f the total number of possible terms is doubled in comparison with the number when the atoms are different. For a reflection with respect to the origin (this being the point bisecting the axis of the molecule) results in an interchange of the states of the two atoms. Sym- metrising or antisymmetrising the wave function of the molecule with respect to an interchange of the states of the atoms, we obtain two terms (with the same A and S), of which one is even and the other odd. Thus we have al- together the same number of even and odd terms. If, on the other hand, both atoms are in the same state, the total number of states is the same as for a molecule with different atoms. An investigation which we shall not give here on account of its length J leads to the following results for the parity of these states. Let N g , N u be the numbers of even and odd terms with given values of A and S. Then if A is odd, N g = N u ; if A is even and S is even (S = 0, 2, 4, ... ), N g = N u + 1 ; if A is even and S is odd (S = 1, 3, 5, ... ), N u = N g +1. Finally, we must distinguish, among the 2 terms, between 2+ and 2 - . Here, if S is even, N g + = iW+l = L+l ; if S is odd, N u + = N g ~+1 = L+l, where L x = L 2 = L. All the 2+ terms are of parity ( — 1) & ', and all 2 - terms are of parity ( — l) s+1 . Besides the problem that we have examined of the relation between the molecular terms and those of the atoms obtained as r -> oo, we may also propose the question of the relation between the molecular terms and those of the "composite atom" obtained as r -> 0, i.e. when both nuclei are brought to a single point (for example, between the terms of the H 2 molecule and those of the He atom). The following rules can be deduced without difficulty. t In particular, we may be discussing the combination of a neutral and an ionised atom. % It can be found in the original paper by E. Wigner and E. Witmer, Zeitschrift fur Physik 51, 859, 1928. 286 The Diatomic Molecule §81 From a term of the "composite" atom having spin S, orbital angular momen- tum L and parity /, we can obtain, on "moving the constituent atoms apart", molecular terms with spin S and angular momentum about the axis A = 0, 1, ... , L, with one term for each of these values of A. The parity of the molecular term is the same as the parity / of the atomic term (g for / = +1 and u for / = — 1). The molecular term with A = is a 2+ term if (-1) L I= + 1, and a 2- term if (-1)^7= -1. PROBLEMS Problem 1. Determine the possible terms for the molecules H 2 , N 2 , 2 , Cl 2 which can be obtained by combining atoms in the normal state. Solution. According to the rules given above, we find the following possible terms: H 2 molecule (atoms in the 2 S state) : *2+ 3 2+ • N 2 molecule (atoms in the *S state) : 1 2+ 32+ 5 2+ 7 2+ • Cl 2 molecule (atoms in the 2 P state) : 2 X E+ X T.- iTT iTT iA„ 2«2+ tt , 3 S - fl> 3n „ 3 n„, 3 A U O a molecule (atoms in the 3 P state) : 2*2+,, *2- M , m g , in., iA ff , 232\, 32-,, m w 3 n„ 3 A, 2 5 S+„, *2- tt , *n g , ^n M , *\. The figures in front of the symbols indicate the number of terms of the type concerned, if this number exceeds unity. Problem 2. The same as Problem 1, but for the molecules HC1, CO. Solution. When unlike atoms are combined, the parity of their states is important also. From formula (31.5) we find that the normal states of the H, O and C atoms are even, while that of the CI atom is odd (see Table 3 for the electron configurations of these atoms). From the rules given above, we have HC1 molecule (atoms in the 2 S g and 2 P U states) : 1,3 S + 1.3TJ; CO molecule (both atoms in the *P g state) : 2 1.3,5 S + 1,3.5 S - 2L3.5II, WA. §81. Valency The property of atoms of combining with one another to form molecules is described by means of the concept of valency. To each atom we ascribe a definite valency, and when atoms combine their valencies must be mutually satisfied, i.e. to each valency bond of an atom there must correspond a valency bond of another atom. For example, in the methane molecule CH4, the four valency bonds of the quadrivalent carbon atom are satisfied by the four univalent hydrogen atoms. In going on to give a physical inter- pretation of valency, we shall begin with the simplest example, the com- bination of two hydrogen atoms to form the molecule H 2 . §81 Valency 287 Let us consider two hydrogen atoms in the ground state ( 2 S). When they approach, the resulting system may be in the molecular state 1 S+ S or 3 E+„. The singlet term corresponds to an antisymmetrical spin wave function, and the triplet term to a symmetrical function. The co-ordinate wave function, on the other hand, is symmetrical for the *E term and antisymmetrical for the 3 E term. It is evident that the ground term of the H 2 molecule can only be the X E term. |For an antisymmetrical wave function (f>(r ly r 2 ) (where r x and r 2 are the radius vectors of the two electrons) always has nodes (since it vanishes for r x = r 2 ), and hence cannot belong to the lowest state of the system. A numerical calculation shows that the electron term X S in fact has a deep minimum corresponding to the formation of a stable H 2 molecule. In the 3 E state, the energy U(r) decreases monotonically as the distance between the nuclei increases, corresponding to the mutual repulsion of the two hydro- gen atomsf (Fig. 29). \U{r) Fig. 29 Thus, in the ground state, the total spin of the hydrogen molecule is zero, 5 = 0. It is found that the molecules of practically all chemically stable compounds of elements of the principal groups have this property. Among inorganic molecules, exceptions are formed by the diatomic molecules 2 (ground state 3 E) and NO (ground state 2 II) and the triatomic molecules N0 2 , C10 2 (total spin S = \). Elements of the intermediate groups have special properties which we shall discuss below, after studying the valency properties of the elements of the principal groups. The property of atoms of combining with one another is thus related to their spin (W. Heitler and H. London 1927). The combination occurs in t Here we ignore the van der Waals attraction forces between the atoms (see §89). The existence of these forces causes a minimum (at a greater distance) on the U(r) curve for the 3 S term also. This minimum, however, is very shallow in comparison with that on the X E curve, and would not be per- ceptible on the scale of Fig. 29. 288 The Diatomic Molecule §81 such a way that the spins of the atoms compensate one another. As a quanti- tative characteristic of the mutual combining powers of atoms, it is convenient to use an integer, twice the spin of the atom. This is equal to the chemical valency of the atom. Here it must be borne in mind that the same atom may have different valencies according to the state it is in. Let us examine, from this point of view, the elements of the principal groups in the periodic system. The elements of the first group (the first column in Table 3, the group of alkali metals) have a spin S = I in the normal state, and accordingly their valencies are unity. An excited state with a higher spin can be attained only by exciting an electron from a completed shell. Accordingly, these states are so high that the excited atom cannot form a stable molecule, f The atoms of elements in the second group (the second column in Table 3, the group of alkaline-earth metals) have a spin S = in the normal state. Hence these atoms cannot enter into chemical compounds in the normal state. However, comparatively close to the ground state there is an excited state having a configuration sp instead of s 2 in the incomplete shell, and a total spin S = 1. The valency of an atom in this state is 2, and this is the principal valency of the elements in the second group. The elements of the third group have an electron configuration s 2 p in the normal state, with a spin S — \. However, by exciting an electron from the completed s-shell, an excited state is obtained having a configuration sp 2 and a spin S = 3/2, and this state lies close to the normal one. Accordingly, the elements of this group are both univalent and tervalent. The first two ele- ments in the group (boron, aluminium) behave only as tervalent elements. The tendency to exhibit a valency 1 increases with the atomic number, and thallium behaves equally as a univalent and as a tervalent element (for example, in the compounds T1C1 and T1C1 3 ). This is due to the fact that, in the first few elements, the binding energy in the tervalent compounds is greater than for the univalent compounds, and this difference exceeds the excitation energy of the atom. In the elements of the fourth group, the ground state has the configuration s 2 p 2 with a spin of 1, and the adjacent excited state has a configuration sp 3 with a spin 2. The valencies 2 and 4 correspond to these states. As in the third group, the first two elements (carbon, silicon) exhibit mainly the higher valency (though the compound CO, for example, forms an exception), and the tendency to exhibit the lower valency increases with the atomic number. In the atoms of the elements of the fifth group, the ground state has the configuration s 2 p 3 with a spin S =3/2, so that the corresponding valency is three. An excited state of higher spin can be obtained only by the transi- tion of one of the electrons into the shell with the next higher value of the principal quantum number. The nearest such state has the configuration sph' and a spin S = 5/2 (by s' we conventionally denote here an s state of an t See the end of this section for the elements copper, silver and gold. §81 Valency 289 electron with a principal quantum number one greater than in the state s). Although the excitation energy of this state is comparatively high, the excited atom can still form a stable compound. Accordingly, the elements of the fifth group behave as both tervalent and quinquevalent elements (thus, nitrogen is tervalent in NH 3 and quinquevalent in HN0 3 ). In the sixth group of elements, the spin is 1 in the ground state (configura- tion s^p 9 ), so that the atom is bivalent. The excitation of one of the p electrons leads to a state s*p 3 s' of spin 2, while the excitation of an s electron in addition gives a state sph'p' of spin 3. In both excited states the atom can enter into stable molecules, and accordingly exhibits valencies of 4 and 6. The first element of the sixth group (oxygen) shows only the valency 2, while the sub- sequent elements show higher valencies also (thus, sulphur in H 2 S, S0 2 , S0 3 is respectively bivalent, quadrivalent and sexivalent). In the seventh group (the halogen group), the atoms are univalent in the ground state (configuration s 2 ^ 5 , spin S = . $). They can, however, enter into stable compounds when they are in excited states having configurations s 2 p*s', s 2 p 3 s'p', sp s s'p' 2 with spins 3/2, 5/2, 7/2 and valencies 3, 5, 7 respec- tively. The first element in the group (fluorine) is always univalent, but the subsequent elements also exhibit the higher valencies (thus, chlorine in HC1, HC10 2 , HC10 3 , HC10 4 is respectively univalent, tervalent, quinquevalent and septivalent). Finally, the atoms of the elements in the group of inert gases have com- pletely filled shells in their ground states (so that the spin S = 0), and their excitation energies are high. Accordingly, the valency is zero, and these elements are chemically inactive.f The following general remark should be made concerning all these discus- sions. The assertion that an atom enters into a molecule with a valency per- taining to an excited state does not mean that, on moving the atoms apart to large distances, we necessarily obtain an excited atom. It means only that the distribution of the electron density in the molecule is such that, near the nucleus of the atom in question, it is close to that in the isolated and excited atom; but the limit to which the electron distribution tends as the distance between the nuclei is increased may correspond to non-excited atoms. When atoms combine to form a molecule, the completed electron shells in the atoms are not much changed. The distribution of the electron density in the incomplete shells, on the other hand, may be considerably altered. In the most clearly denned cases of what is called heteropolar binding, all the valency electrons pass over from their own atoms to other atoms, so that we t The elements xenon and radon (and less easily krypton) nevertheless form stable compounds with fluorine. These valencies may be due to a transfer of electrons from the outermost complete shell to the incomplete / or d states, whose energies are comparatively near. There is also an attraction which occurs in the interaction of an inert gas atom with an excited atom of the same element. This is due to the doubling in the number of possible states obtained on bringing together two atoms, if these atoms are of the same element but in different states (see §80). The transi- tion of the excitation from one atom to the other here replaces the exchange interaction which brings about the ordinary valency. The molecule He2 is an example of such a molecule. The same type of bond occurs in molecular ions composed of two similar atoms (for instance, H2 + ). 290 The Diatomic Molecule §81 may say that the molecule consists of ions with charges equal (in units of e) to the valency. The elements of the first group are electropositive : in hetero- polar compounds they lose electrons, forming positive ions. As we pass to the subsequent groups the electropositive character of the elements becomes gradually less marked and changes into electronegative character, which is present to the greatest extent in the elements of the seventh group. Regard- ing heteropolarity the same remark should be made as was made above con- cerning excited atoms in the molecule. If a molecule is heteropolar, this does not mean that, on moving the atoms apart, we necessarily obtain two ions. Thus, from the molecule KC1 we should in fact obtain the ions K+ and CI - , but the molecule NaCl gives in the limit the neutral atoms Na and CI (since the affinity of chlorine for an electron is greater than the ionisation potential of potassium but less than that of sodium). In the opposite limiting case of what is called homopolar binding, the atoms in the molecule remain neutral on the average. Homopolar molecules, un- like heteropolar ones, have no appreciable dipole moment. The difference between the heteropolar and homopolar types is purely quantitative, and any intermediate case may occur. Let us now turn to the elements of the intermediate groups. Those of the palladium and platinum groups are very similar to the elements of the principal groups as regards their valency properties. The only difference is that, owing to the comparatively deep position of the d electrons inside the atom, they interact only slightly with the other atoms in the molecule. As a result, "unsaturated" compounds, whose molecules have non-zero spin (though in practice not exceeding £), are often found among the compounds of these elements. Each of the elements can exhibit various valencies, and these may differ by unity, and not only by two as with the elements of the principal groups (where the change in valency is due to the excitation of some electron whose spin is compensated, so that the spins of two electrons are simultaneously released). The elements of the rare-earth group are characterised by the presence of an incomplete / shell. The / electrons lie much deeper than the d electrons, and therefore take no part in the valency. Thus the valency of the rare- earth elements is determined only by the s and p electrons in the incomplete shells.f However, it must be borne in mind that, when the atom is excited, / electrons may pass into s and p states, thereby increasing the valency by one. Hence the rare-earth elements too exhibit valencies differing by unity (in practice they are all tervalent and quadrivalent). The elements of the actinium group occupy a unique position. Actinium and thorium have no / electrons, and their valencies involve d electrons. In their chemical properties they are therefore analogous to elements of the palladium and platinum groups, not to the rare earths. The uranium atom in t The d electrons which are found in the incomplete shells of the atoms of some rare-earth elements are unimportant, since these atoms in practice always form compounds in excited states where there are no d electrons. §81 Valency 291 the normal state contains / electrons, but in its compounds it too has no / electrons. Finally, the atoms of the elements neptunium, plutonium, americium and curium contain/ electrons in compounds also, but the electrons which participate in their valencies are again s and d electrons. In this sense they are homologues of uranium. The maximum possible number of "un- paired" s and d electrons is one and five respectively, and so the maximum valency of elements in the actinium group is six, whereas the maximum valency of the rare-earth elements (with 5 and p electrons participating in the valency) is 1 + 3 = 4. The elements of the iron group occupy, as regards their valency properties, a position intermediate between the rare-earth elements and those of the palla- dium and platinum groups. In their atoms, the d electrons lie comparatively deep, and in many compounds take no part in the valency bonds. In these compounds, therefore, the elements of the iron group behave like rare-earth elements. Such compounds include those of ionic type (for instance FeCl 2 , FeCl 3 ), in which the metal atom enters as a simple cation. Like the rare- earth elements, the elements of the iron group can show very various valencies in these compounds. Another type of compound of the iron-group elements is formed by what are called complex compounds. These are characterised by the fact that the atom of the intermediate element enters into the molecule not as a simple ion, but as part of a complex ion (for instance the ion Mn0 4 ~~ in KMn0 4 , or the ion Fe(CN) 6 4_ in K 4 Fe(CN) 6 ). In these complex ions, the atoms are closer together than in simple ionic compounds, and in them the d electrons take part in the valency bond. Accordingly, the elements of the iron group behave in complex compounds like those of the palladium and platinum groups. Finally, it must be mentioned that the elements copper, silver and gold, which in §73 we placed among the principal groups, behave as intermediate elements in some of their compounds. These elements can exhibit valencies of more than one, on account of a transition of an electron from a d shell to a p shell of nearly the same energy (for example, from 3d to 4p in copper). In such compounds the atoms have an incomplete d shell, and hence behave as intermediate elements : copper like the elements of the iron group, and silver and gold like those of the palladium and platinum groups. PROBLEM Determine the electron terms of the molecular ion H2 + obtained when a hydrogen atom in the normal state combines with an H + ion, for distances R between the nuclei large compared with the Bohr radius. Solution. This problem is analogous in form to §50, Problem 3: instead of two one-dimensional potential wells we have here two three-dimensional wells (round the two nuclei) with axial symmetry about the line joining the nuclei. The levelf E0 = —i (the ground level of the hydrogen atom) is split into two levels U g (R) and U U (R) (the terms 2 S ff + and 2 E U +), corresponding to the electron wave functions 1 <A<7,uO*y>#) = — -[M x >y> z )±M- x >y> z )l v 2 t Here we are using'atomic units. 292 The Diatomic Molecule §81 which are symmetrical and antisymmetrical about the plane x = which bisects the line joining the nuclei (which are at (±iR, 0, 0)). Here *f>o(x, y, z) is the wave function of the electron in one of the potential wells. Exactly as in §50, Problem 3, we find ndfo 00 dydz, (1) 8x where the integration is over the plane x = 0. The function ^o (corresponding to motion around nucleus 1, say, at x = %R) is sought in the form h = -^e-r*, (2) V 77- where a is a slowly varying function (for a hydrogen atom, a = 1). The function ^o must satisfy Schrodinger's equation / 1 1 1 \ £A<A+(-£-- + - + - <£ = o, \ R n r 2 / (3) where n, r% are the distances of the electron from nuclei 1 and 2. In this equation the total energy of the electron is Eo — l/R, since Eq itself includes the energy 1/R of the Coulomb repulsion of the nuclei. Since the function tfio decreases rapidly away from the #-axis, only the region where y and z are small compared with R is important in the integral (1). For y, z <^ R, substitution of (2) in (3) gives da a a — + - = 0; dx \R+x R here we have neglected the second derivatives of the slowly varying function a and put ?2 S iR+x. The solution of this equation which becomes unity as x -*■ %R (i.e. in the neighbourhood of nucleus 1) is 2R exp R+2x a->} Formula (1) now gives Ug, u -Eo=+— I e-^i . 2ttt 1 dr ± - S ire J R/2 = +(2/e)Re-R At sufficiently large distances this expression decreases exponentially and becomes less than the effect in the second approximation with respect to the dipole interaction of the H atom and the H + ion. Since the polarisability of the hydrogen atom in the normal state is 9/2 (see (77.9)), and the field of the H"' ion is $ = 1/R 2 , the corresponding interaction energy is —9/4R 4 , and when this is taken into account we have U g , u (R) - Eo = + -Re-* - — . (4) e 4/c 4 The second term becomes comparable with the first when R = 10-8. It may also be noted that the term U u has a minimum of — 5 -8 x 10~ 5 atomic unit ( — 1 -6 x 10~ 3 eV) when R = 12-6.f t This minimum, which is due to van der Waals forces, is very shallow compared with that of the term U g (R) which corresponds to the normal state of the stable ion H2 + : the latter minimum is -0-60 atomic unit (-16-3 eV), at R = 2-0. §82 Vibrational and rotational structures of singlet terms 293 §82. Vibrational and rotational structures of singlet terms in the diatomic molecule As has been pointed out at the beginning of this chapter, the great differ- ence in the masses of the nuclei and the electrons makes it possible to divide the problem of determining the energy levels of a molecule into two parts. We first determine the energy levels of the system of electrons, for nuclei at rest, as functions of the distance between the nuclei (the electron terms). We can then consider the motion of the nuclei for a given electron state ; this amounts to regarding the nuclei as particles interacting with one another in accordance with the law U n (r), where U n is the corresponding electron term. The motion of the molecule is composed of its translational displacement as a whole, together with the motion of the nuclei about their centre of mass. The translational motion is, of course, without interest, and we can regard the centre of mass as fixed. For convenience of discussion, let us first consider the electron terms in which the total spin S of the molecule is zero (the singlet terms). The problem of the relative motion of two particles (the nuclei) which interact according to the law U(r) reduces, as we know, to that of the motion of a single particle of mass M (the reduced mass of the two particles) in a centrally symmetric field U(r). By U(r) we mean the energy of the electron term considered. The problem of motion in a centrally symmetric field U(r), however, reduces in turn to that of a one-dimensional motion in a field where the effective energy is equal to the sum of U(r) and the centrifugal energy. We denote by K the total angular momentum of the molecule, composed of the orbital angular momentum L of the electrons and the angular momen- tum of the nuclei. Then the operator of the centrifugal energy of the nuclei B(r)(K-L)\ where we have introduced the notation B(r) = H 2 /2Mr 2 (82.1) for a convenient simplification of the formulae in the theory of diatomic molecules. Averaging this quantity (for a given r), we obtain the centrifugal energy as a function of r, which must appear in the effective potential energy U K (r). Thus U K (r) = t/(r)+5(r)(K-L)*, where the line denotes the average mentioned. Expanding the square and recalling that the square K 2 of a conserved total angular momentum has the definite value K(K+1) (where K is integral), we can rewrite this expression in the form U K (r) = U(r)+B(r)K(K+l)+B(r)(l}-2L . K); (82.2) we omit the line over the quantity K, since it is conserved. 294 The Diatomic Molecule §82 In a state with a definite value of L z = A, the mean values of the other two components of the orbital angular momentum are zero, L x = L y = 0; this follows at once from the fact that, in a representation in which L z is diagonal, the diagonal matrix elements of the operators L x and L are zero (see §27). Hence the mean value of the vector L is plirected along the sr-axis and we can write L = nA, where n is a unit vector along the axis of the molecule. Next, in classical mechanics the angular momentum of a system of two particles (the nuclei) is directed perpendicular to the line joining them ; in quantum mechanics the same is true for the angular momentum operator. Hence we can write (R— L) . n = 0, and R . n = L . n. Hence, for the eigenvalues, K.n=L.n=A. (82.3) Thus the projection of the total angular momentum K on the axis of the mole- cule is also A. Hence it follows that, in a state with a given value of A, the quantum number K can take only values from A upwards : K > A. (82.4) Finally, substituting in (82.2) L . K = An . K = A 2 , we obtain U K (r) = U(r)+B(r)K(K+l)+B(r)(V-2A*). (82.5) The last term on the right-hand side is some function of r, depending only on the electron state, and not on the quantum number K. This function can be included in the energy U(r), and (82.5) then takes the form U K (r) = U(r)+B(r)K(K+l). (82.6) On solving the one-dimensional Schrodinger's equation with this potential energy, we obtain a series of energy levels. We arbitrarily number these levels (for each given K) in order of increasing energy, using a number v = 0, 1, 2, ... ; v = corresponds to the lowest level. Thus the motion of the nuclei causes a splitting of each electron term into a series of levels characterised by the values of the two quantum numbers K and v. The number of these levels (for a given electron term) may be either finite or infinite. If the electron state is such that, as r -> oo, the molecule becomes two isolated neutral atoms, then as r -> oo the potential energy U(r) (and therefore U K (r)) tends to a constant limiting value U(oo) (the sum of the energies of the two isolated atoms) more rapidly than 1/r tends to zero (see §89). The number of levels in such a field is finite (see §18), though in actual molecules it is very large. The levels are so distributed that, for any given value of K, there is a definite number of levels (with different values of v), while the number of levels with the same K diminishes as K increases, until a value of K is reached for which there are no levels at all. §82 Vibrational and rotational structures of singlet terms 295 If, on the other hand, as r -» oo the molecule disintegrates into two ions, at large distances U(r) — U(oo) becomes the energy of the attraction of the ions according to Coulomb's law (^ 1/r). In such a field there is an infinite number of levels, which become closer and closer as we approach the limiting value E/(oo). We may remark that, for the majority of molecules, the previous case is found in the normal state ; only a comparatively small number of mole- cules become pairs of ions when their nuclei are moved apart. The dependence of the energy levels on the quantum numbers cannot be completely calculated in a general form. Such a calculation is possible only for low excited levels which lie not too far above the ground level.f Small values of the quantum numbers K and v correspond to these levels. It is with such levels that we are in fact most often concerned in the study of molecular spectra, and hence they are of particular interest. The motion of the nuclei in slightly excited states can be regarded as small vibrations about the equilibrium position. Accordingly we can expand U(r) in a series of powers of £ = r— r e , where r e is the value of r for which U(r) has a minimum. Since U'(r e ) — 0, we have as far as terms of the second order U(r) = t/ 8 +|Mov^, where U e = U(r e ), and a> e is the frequency of the vibrations. J In the second term in (82.6) — the centrifugal energy — it is sufficient to put r = r e , since it already contains the small quantity K(K+1). Thus we have U K (r) = U.+BJC{K+l)+tM<**?, (82.7) where B e = W>\2Mr? = h % \71 is what is called the rotational constant (I ='"Mr e 2 is the moment of inertia of the molecule). The first two terms in (82.7) are constants, while the third corresponds to a one-dimensional harmonic oscillator. Hence we can at once write down the required energy levels : E = U e +B e K(K+l)+hco e (v+l). (82.8) Thus, in the approximation considered, the energy levels are composed of three independent parts : E = E el +E r +E v . (82.9) Here E a = U e is the electron energy (including the energy of the Coulomb interaction of the nuclei for r = r e ), E r =B e K{K+\) (82.10) t We refer always to levels belonging to the same electron term. J We here use the notation customary in the theory of diatomic molecules. 296 The Diatomic Molecule §82 is the rotational energy from the rotation of the molecule, f and E v = h<* e {v+\) (82.11) is the energy of the vibrations of the nuclei within the molecule. The number v denumerates, by definition, the levels with a given K in order of increasing energy; it is called the vibrational quantum number. For a given form of the potential energy curve U(r), the frequency <o e is inversely proportional to ^JM. Hence the intervals AE V between the vibrational levels are proportional to 1[\/M. The intervals AE r between the rotational levels contain in the denominator the moment of inertia /, and are therefore proportional to \jM. The intervals AE el between the electron levels, however, are independent of M, like the levels themselves. Since mjM (m being the electron mass) is a small parameter in the theory of diatomic molecules, we see that AE el > AE V > AE r . Thus the distribution of the energy levels of the molecule is rather unusual. The vibrational motion of the nuclei splits the electron terms into levels lying comparatively close together. These levels, in turn, exhibit a fine splitting due to the rotational motion of the molecule. J In subsequent approximations, the separation of the energy into indepen- dent vibrational and rotational parts is impossible; rotational- vibrational terms appear, which contain both K and v. On calculating the successive approximations, we should obtain the levels E as an expansion in powers of the quantum numbers K and v. We shall calculate here the next approximation after (82.8). To do this, we must continue the expansion of U(r) in powers of £ up to terms of the fourth order (cf. the problem of an anharmonic oscillator in §38). Similarly, the expansion of the centrifugal energy is extended as far as the terms in | 2 . We then obtain U K {r) = U e +iM^e+(h 2 l2Mr e ^)K(K+l)~ -a£*+b?-(h 2 IMr*)K(K+ l)£+(3k*l2Mr t *)K(K+ 1)£ 2 . (82.12) Let us now calculate the correction to the eigenvalues (82.8), using perturbation theory and regarding the last four terms in (82.12) as the per- turbation operator. Here it is sufficient, for the terms in | 2 and £ 4 , to take the first approximation of perturbation theory, but for those in f and £ 3 we must t A rotating system of two rigidly connected particles is often called a rotator. Formula (82.10) gives the quantum-mechanical energy levels for a rotator. The wave functions of the stationary states of a rotator evidently correspond to the case A = and are ordinary spherical harmonic functions (see the Problem at the end of this section). J As an example, we give the values of U e , hu>e and B e (in electron-volts) for a few molecules : H 2 N 2 2 -U e 4-7 7-5 5-2 Hcoe 0-54 0-29 0-20 10 3 x5 e 7-6 0-25 0-18 §82 Vibrational and rotational structures of singlet terms 297 calculate the second approximation, since the diagonal matrix elements of | and | 3 vanish identically. All the matrix elements needed for the calculation are derived in §23 and in §38, Problem 3. As a result, we obtain an expres- sion which is usually written in the form E - E tl +ha>Jiv+\)-x-hco 6 {v+}tf+B v K{K+ \)-D € K%K+ 1) 2 , (82.13) where B v = B e — a. e (v+%) = B —cc e v. (82.14) The constants x ey B e , <x e , D e are related to the constants appearing in (82.12) by B e = /* 2 /27, D = AB e z lh*<* e \ 6B*Y ah I 2 \ 3 /h Yf 5 a% b\ (82.15) The terms independent of v and K are included in E el . PROBLEM Determine the angular part of the wave function for a diatomic molecule with zero spin (F. Reiche 1926). Solution. The required functions are just the eigenfunctions of the total angular momen- tum K of the molecule. The operator of the total angular momentum is the sum R =rxp+ Sr a xp , where p is the linear momentum of the relative motion of the nuclei, r the radius vector be- tween them, r„ and p tt the radius vectors and linear momenta of the electrons (relative to the centre of mass of the molecule). Introducing the polar angle 6 and the azimuthal angle <f> of the axis of the molecule relative to a fixed system of co-ordinates *, y, z, we have for the components of the operator & expressions similar to (26.14), (26.15), so that & = e x H \-i cot0 — )+L+, — + i cot e— )+£_, dd d<f>J (i) it. ■ d ' t ■i \-L z , where L . = Lx + ilj. . / 8 8 \ are the operators of the angular momenta of the electrons; the primes on djdd and djd<j> signify that the differentiation is to be performed for constant x a , ya, Za- Besides the fixed system of co-ordinates x, y, z, we introduce a moving system £, t), £, with the same origin, the £-axis directed along the axis of the molecule, and the £-axis lying in the 298 The Diatomic Molecule §82 «y-plane. The co-ordinates £ a , ya, U of the electrons in this system are related to the co- ordinates X a , ya, Za by € a = —x a sin <f>+y a coscf), -q a = — x a cos 6 cos<f>— y a cos0 sin^+.s'a sin0, t a = x a sin 6 cos <f>+y a sin 9 sin (f>+z a cos 6. Using these formulae, we can transform the derivatives: 8 8 d = sin0 [-cosd , etc., fea fya %a d 8 Z\ 8 d 8$ a 86 8 Va 86 8tJ 8 \? / d 8 \ where 9/90 and djd<f> (unprimed) denote differentiation for constant £ a , 17a, U- As a result, we have for the operators of the components of the total angular momentum relative to the fixed system the expressions K+ = e«P [ — \-i cot0 — )+ L,, \86 8<f>J sin 6 ^ £ z = -i8j8<j>, where is the operator of the angular momentum of the electrons about the axis of the molecule. Let *PnAKM K = <f>nAK(€ a > Va, £a"> r)p n AK(r)@ A KM K (9)^>M K (<f>) (3) be the wave function of a state with definite values of the absolute value K and ^-component Mk of the total angular momentum of the molecule, a nd a definite value A of the ^-component of the electron angular momentum ; n denotes the ass embly of the remaining quantum num- bers which determine the state of the molecule. <f>nAK is the electron wave function, depend- ing on r as a parameter, p n AK is the "radial part" of the nuclear wave function, &AKMK is the required function of the angle 0, and the dependence of ifi on the angle <f> is obvious : 1 *>**&) = —^rr eiMR<l> ' v( 27r ) When the operators J2" z , £j act on the function (3), we can replace them by their eigenvalues §83 Multiplet terms 299 Mr, A, so that /d \ e x< f g = e m M K cot0 ) + A, \dd / sin 9 ^ ,/ 3 \ e"^ it- = e-^( M x cot0 )+ A. \ 36 J sin9 The subsequent argument exactly follows that at the end of §28. When the operator &+ acts on the function «A„Akk (with Mr = K), the result is zero; hence we have the equation d A \ Kcot6+ )®AKK = 0, G whose solution ist (2*1+1) 2*k+i(K+A)\(K-A)\ the function is normalised by the condition ®AKK=(-i) K I (2 ^ +1)! x(l-cos^- A )/2(l + cos^+ A )/2; (4) AKK K ' V 2^+UK+A)\(K-A)\ je AKK 2 sin ede = 1, and the normalisation integral reduces to Euler's beta function. The remaining functions are then calculated from the formula - — — -Qakm =R- k - Mk &akk, V (K+M K )\ and as a result we obtain / (2K+1)\(K+M K )\ (I - cos 8)( a -Mk) 12 G\ KT71l/r = t-.i\K / 1 _ 1 x X VAKM K \ ) ,J ( K+fi)] ( K _ A) l( K _ MK )\ (l + cOsdy A +M«)V KQ k K-Mk -i ) (l-cos^- A (l+cos^+ A . d cos 9 J J For A = these functions become ordinary spherical harmonic functions, as they should: ®0KM = constant x P^*(cos 6), and are the wave functions of a rotator (eigenfunctions of the free angular momentum K). §83. Multiplet terms. Case a Let us now turn to the question of the classification of molecular levels with non-zero spin S. In the zero-order approximation, when relativistic effects are entirely neglected, the energy of the molecule, like that of any system of particles, is independent of the direction of the spin (the spin is "free"), and this results in a (2S+l)-fold degeneracy of the levels. When relativistic f The choice of the phase factor accords with the definition of the eigenfunctions of the free angular momentum (§28), which are obtained for A = 0. 300 The Diatomic Molecule §83 effects are taken into account, however, the degenerate levels are split, and the energy consequently becomes a function of the projection of the spin on the axis of the molecule. We shall refer to relativistic interactions in mole- cules as the spin-axis interaction. The chief part in this is played (as in the case of atoms) by the interaction of the spins with the orbital motion of the electrons.^ The nature and classification of molecular levels depend markedly on the relative parts played by the interaction of the spin with the orbital motion, on the one hand, and the rotation of the molecule, on the other. The part played by the latter is characterised by the distances between adjacent rota- tional levels. Accordingly, we have to consider two limiting cases. In one, the energy of the spin-axis interaction is large compared with the energy differences between the rotational levels, while in the other it is small. The first case is usually called case (or coupling type) a, following Hund, and the second is called case b. Case a is the one most often found. An exception is formed by the 2 terms, where case b chiefly occurs, since the effect of the spin-axis interaction is very small for these terms! (see below). For other terms, case b is some- times found in the lightest molecules, since the spin-axis interaction is here comparatively weak, while the distances between the rotational levels are large (the moment of inertia being small). Of course, cases intermediate between a and b are also possible. It must also be borne in mind that the same electron state may pass continuously from case a to case b as the rotational quantum number changes. This is due to the fact that the distances between adjacent rotational levels increase with the rotational quantum number, and hence, when this is large, the distances may become large compared with the energy of the spin-axis coupling (case b), even if case a is found for the lower rotational levels. In case a, the classification of the levels is in principle little different from that of the terms with zero spin. We first consider the electron terms for nuclei at rest, i.e. we neglect rotation entirely; besides the projection A of the orbital angular momentum of the electrons, we must now take into account the projection of the total spin on the axis of the molecule. This projection is denoted by|| S; it takes the values S, 5-1, ... , -S. We arbitrarily regard 2 as positive when the projection of the spin is in the same direction as that of the orbital angular momentum about the axis (we recall that A denotes the absolute value of the latter). The quantities A and S combine to give the total angular momentum of the electrons about the axis of the molecule : Q=A+S; (83.1) f Besides the spin-orbit and spin-spin interactions there is also an interaction of the spin and orbital motion of the electrons with the rotation of the molecule. This part of the interaction is very small, however, and it is of possible interest only for terms with spin S = J (see §84). X A special case is the normal electron term of the molecule 2 (the term S S). For this we have a type of coupling intermediate between a and b (see §84, Problem 3). || Not to be confused with the symbol for terms with A == 0. §83 Multiplet terms 301 this takes the values A+S, A+S-l, ... , A-S. Thus the electron term with orbital angular momentum A is split into 25+1 terms with different values of ft; this splitting, as with atomic terms, is called the^ine structure or multiplet splitting of the electron levels. The value of H is usually indicated as a suffix to the symbol for the term: thus, for A = 1, S = \ we obtain the terms 2 IIi/2, 2 n 3/2 . When the motion of the nuclei is taken into account, vibrational and rota- tional structures appear in each of these terms. The various rotational levels are characterised by the values of the quantum number /, which gives the total angular momentum of the molecule, including the orbital and spin angular momenta of the electrons and the angular momentum of the rotation of the nuclei.f This number takes all integral values from \€i\ upwards: 7>|0|, ( 83 - 2 ) which is an obvious generalisation of (82.4). Let us now derive quantitative formulae to determine the molecular levels in case a. First of all, we consider the fine structure of an electron term. In discussing the fine structure of atomic terms in §72, we used formula (72.4), according to which the mean value of the spin-orbit interaction is proportional to the projection of the total spin of the atom on the orbital angular momentum vector. Similarly, the spin-axis interaction in a diatomic molecule (averaged over electron states for a given distance r between the nuclei) is proportional to the projection S of the total spin of the molecule on its axis, so that we can write the split electron term in the form C/(r)+^(r)S, where U(r) is the energy of the original (unsplit) term, and A(r) is some func- tion of r; this function depends on the original term (and in particular on A), but not on S. Since one usually uses the quantum number Q. and not 2, it is more convenient to put AQ. in place of AZ; these expressions differ by AA, which can be included in U(r). Thus we have for an electron term the expression /M „ U(r)+A(r)Cl. (83.3) We may notice that the components of the split term are equidistant from one another: the distance between adjacent components (with values of Q. differing by unity) is A(r), independent of £1. It is easy to see from general considerations that the value of A for S terms is zero. To show this, we perform the operation of changing the sign of the time. The energy must then remain unchanged, but the state of the mole- cule changes in that the direction of the orbital and spin angular momenta about the axis is reversed. In the energy ^4(r)S, the sign of S is changed, and if the energy remains unchanged A(r) must change sign. If A # 0, we can f The notation K is, as usual, reserved for the total angular momentum of the molecule without allowance for its spin. In case a there is no quantum number K, since the angular momentum K is not even approximately conserved. 302 The Diatomic Molecule §83 draw no conclusions regarding the value of A(r), since this depends on the orbital angular momentum, which itself changes sign. If A = 0, however, we can say that A(r) is certainly unchanged, and consequently it must vanish identically. Thus, for the S terms, the spin-orbit interaction causes no split- ting in the first approximation; splitting (proportional to S 2 ) would occur only on taking account of this interaction in the second approximation or the spin-spin interaction in the first approximation, and would be relatively small. This is the reason for the fact, already mentioned, that case b usually occurs for 21 terms. When the multiplet splitting has been determined, we can take account of the rotation of the molecule as a perturbation, just as in the derivation given at the beginning of §82. The angular momentum of the rotation of the nuclei is obtained from the total angular momentum by subtracting the orbital angular momentum and spin of the electrons. Hence the operator of the centrifugal energy now has the form B(r)(J-L-S)\ Averaging this quantity with respect to the electron state and adding to (83.3), we obtain the required effective potential energy Uj(r): Uj(r) = U(r)+A(r)Q+B(r)(]-L-Sf = tf(r)+^(r)Q+£(r)[J 2 -2J.(L+S)+L 2 +2L. S+S 2 ]. The eigenvalue of J 2 is /(/ + 1). Next, by the same argument as in §82, we have L = nA, S = nS, (83.4) and also (J— L— S) . n = 0, whence we have for the eigenvalues J.n =(L+S).n =A+S = O. (83.5) Substituting these values, we find Uj(r) = U(r)+A(r)£l+B(r)[J(J+l)-2Q. 2 +V+2L.S+S 2 ]. The averaging with respect to the electron state is effected by means of the wave functions of the zero-orderf approximation. In this approximation, however, the magnitude of the spin is conserved, and hence S 2 = S^S+l). The wave function is the product of the spin and co-ordinate functions ; hence the averaging of the angular momenta L and S takes place independently, and we obtain L.S =An.S =AS. f That is, the zero-order approximation with respect to both the effect of the rotation of the mole- cule and the spin-axis interaction. §84 Multiplet terms 303 Finally, the mean value of the squared orbital angular momentum L 2 is independent of the spin, and is some function of r characterising the given (unsplit) electron term. All the terms which are functions of r but indepen- dent of J and S can be included in U(r), while the term proportional to S (or, what is the same thing, to Q.) can be included in the expression A{r)Cl. Thus we have for the effective potential energy the formula Uj(r) = U(r)+A(r)£l+B(r)[J(J+l)-2&]. (83.6) The energy levels of the molecule can be obtained from this by the same method as in §82 when using the formula (82.6). Expanding U(r) and A(r) in series of powers of £, and retaining the terms up to and including the second order in the expansion of U(r), but only the terms of zero order in the second and third terms, we obtain the energy levels in the form E = U 6 +A 6 a+h<* e {v+\)+B e [](J+\)-2Wl (83.7) where A e = A(r e ) and B e are constants characterising the given (unsplit) electron term. On continuing the expansion to higher terms, we obtain a series of terms in higher powers of the quantum numbers, but we shall not pause to write these out here. §84. Multiplet terms. Case b Let us now turn to case b. Here the effect of the rotation of the molecule predominates over the multiplet splitting. Hence we must first consider the effect of rotation, neglecting the spin-axis interaction, and then the latter must be taken into account as a perturbation. In a molecule with "free" spin, not only the total angular momentum J but also the sum K of the orbital angular momentum of the electrons and the angular momentum of the nuclei are conserved; the latter is related to J by J=K+S. (84.1) The quantum number K distinguishes different states of a rotating molecule with free spin that are obtained from a given electron term. The effective potential energy Uj^r) in a state with a given value of K is evidently deter- mined by the same formula (82.6) as for terms with S = 0: U K (r) = U(r)+B(r)K(K+l), (84.2) where K takes the values A, A+l, ... . When the spin-axis interaction is included, there is a splitting of each term into 2S+1 terms in general (or 2K+1 if K < S), which differ in the value of the total angular momentum! /. According to the general rule for the addition of angular momenta, the number / takes (for a given K) values f In case b, the projection n.S of the spin on the axis of the molecule does not have definite values, so that there is no quantum number S (or Q). 304 The Diatomic Molecule §84 from K+S to \K-S\: \K-S\<J<K+S. (84.3) To calculate the energy of the splitting (in the first approximation of per- turbation theory), we must determine the mean value of the operator of the spin-axis interaction energy for the state in the zero-order approximation (with respect to this interaction). In the case considered, this means averag- ing with respect to both the electron state and the rotation of the molecule (for a given r). The result of the first averaging is, as we know, an operator of the form A(r)n . S, which is proportional to the projection n . S of the spin operator on the axis of the molecule. Next we average this operator with respect to the rotation of the molecule, taking the direction of the spin vector to be arbitrary; then n . § = h . S. The mean value n is a vector which, from considerations of symmetry, must have the same direction as the "vector" R, the only vector which characterises the rotation of the molecule. Thus we can write n = constant x £. The coefficient of proportionality is easily determined by multiplying both sides of this equation by R; noting that the eigenvalues of n . K and K 2 are respectively A (see (82.3)) and K(K+ 1), we find the constant to be AIKIK+ 1). Thus n.S =AR.S/K(K+l). Finally, the eigenvalue of the product K . S, according to the general formula (31.2), is K. S = W(J+V-K(K+1)-S(S+1)]. (84.4) As a result, we arrive at the following expression for the required mean value of the energy of the spin-axis interaction : Ar)A[J(J+ 1)-S(S+ \)-K{K+ 1)]I2K(K+ 1) = Ar)A[(J-S)(J+S+l)]l2K(K+l)-lA(r)A. This expression must be added to the energy (84.2). The term \A(r)A, being independent of K and /, can be included in U(r), so that we have finally for the effective potential energy the expression U K (r) = U(r)+B(r)K(K+l)+A(r)A(J-S)(J+S+l)l2K(K+l). (84.5) An expansion in powers of £ = r— r e gives, in the usual manner, an expres- sion for the energy levels of the molecule in case b : E = U e +h<* e {v+\)+B e K{K+\)+A e A{J-S){J+S+\)}K{K+\). (84.6) §84 Multiplet terms 305 As has been pointed out in the previous section, the spin-orbit interaction for 2 terms does not give a multiplet splitting in the first approximation, and to determine the fine structure we must take into account the spin-spin interaction, whose operator is quadratic with respect to the spins of the elec- trons. We are at present interested not in this operator itself, but in the result of averaging it with respect to the electron state of the molecule, as was done for the operator of the spin-orbit interaction. It is evident from considera- tions of symmetry that tie required averaged operator must be proportional to the squared projection of the total spin of the molecule on the axis, i.e. it can be written in the form a(r)(S.n) 2 , (84.7) where a(r) is again some function of the distance r, characterising the given electron state. Symmetry allows also a term proportional to S 2 , but this is immaterial since the absolute value of the spin is just a constant. We shall not pause here to derive the lengthy general formula for the splitting due to the operator (84.7) ; in Problem 1 of this section we give the derivation of the formula for triplet 21 terms. The doublet S terms form a special case. According to Kramers' theorem (§60), the double degeneracy in a system of particles with total spin S = \ certainly persists, even when the internal relativistic interactions in the system are fully allowed for. Hence the 2 I! terms remain unsplit, even when we take account of both the spin-orbit and the spin-spin interaction, and in any approximation. The splitting is obtained here only by taking into account the relativistic interaction of the spin with the rotation of the molecule; this effect is very small. The averaged operator of this interaction must evidently be of the form y& . S, and its eigenvalues are determined by the formula (84.4), in which we must put S = $, / = K±%. As a result, we obtain for the 2 S terms the formula E = U e +h<* e (v+$+B e K(K+l)±b>(K+$; (84.8) a constant — £y is included in U e . PROBLEMS Problem 1. Determine the multiplet splitting of a 3 2 term in case b (H. A. Kramers 1929). Solution. The required splitting is determined by the operator (84.7), which must be averaged with respect to the rotation of the molecule. We write it in the form <x e ninicSiSk, where oe e = oc(ro). Since the vector S is conserved, only the products mmc need be averaged. According to the formula derived in §29, Problem 2, we have KiKjc + KjcKi ntnjc = h ... ; (2a:-i)(2*:+3) here we have not written out the terms proportional to 8a whose contribution to the energy is independent of J and therefore does not cause any splitting of the type under consideration. 306 The Diatomic Molecule §84 Thus the splitting is given by the operator (2K-l)(2K+3) Since S commutes with &, SiSicRtK* = SiKiSjcK* = (S . K)2, where the eigenvalue S . K is given by (84.4). We also have StSjcKjcKi = SiSjcKiKjc+iSiSjceMiKi = (S . E)*+U£ i S t -£ k S t )ie ai &i = (S . K) 2 +^en c ien, ;m i§ m Ki = (S.K)2 + S.K. The values/ = K, K± 1 correspond to the three components E K of the triplet 3 S (S = 1). For the intervals between these components we find K+\ K Ek+i - E K = — g< £#_! ~ E K = —a. '2K+1' e 2K-l Problem 2. Determine the energy of a doublet term (with A#0) for cases intermediate between a and 6 (E. Hill and J. H. van Vleck 1928). Solution. Since the rotational energy and the energy of the spin-axis interaction are supposed of the same order of magnitude, they must be considered together in perturbation theory, so that the perturbation operator is of the formf As wave functions in the zero-order approximation it is convenient to use those of states in which the angular momenta .Kand / have definite values (i.e. those of case b). Since S — £ for a doublet term, the quantum number K, for a given /, can take the values K = /±J. To construct the secular equation, we must calculate the matrix elements VZIk v (« denoting the assembly of quantum numbers defining the electron term), where K,K' take the above values. The matrix of the operator K 2 is diagonal ; the diagonal elements are K(K+ 1). The matrix elements of n . S are calculated from the general formula (109.5), in which we must putji = S, 72 = K; the matrix elements of n are given by (87.1). Calculation gives the secular equation B e (J+m+%)-AAI(2J+ 1)-£W AvUm) 2 -A 2 ]/(2/+ 1) ^V[C/+i) 2 -A 2 ]/(2/+l) B e (J+W-l)+AAI(2J+l)-EV Solving this equation and adding E^ to the unperturbed energy, we have E = u e +h^v+±)+B e j{j+\)±^[B*u+W-ABA+lA 2 Y, a constant \B e is included in U e . The inequality A e ^> B e J corresponds to case a, and the opposite one to case b. Problem 3. Determine the intervals between the components of a triplet level 3 Sina case intermediate between a and b. = o. t The averaging with respect to vibrations must be done before that with respect to rotation. Hence, restricting ourselves to the first terms of the expansions in £, we have replaced the functions B(r ) and A(r) by the values B e and A e , and the unperturbed energy levels are 2?(°> = U e + Hco e (v + J). §85 Multiplet terms 307 Solution. As in Problem 2, the rotational energy and the energy of the spin-spin inter- action are considered together in the perturbation theory. The perturbation operator is of the form t = 5 e & 2 +a 6 (n.S) 2 . As wave functions in the zero-order approximation we use those of case b. The matrix elements (n . S)f . (we omit all suffixes with respect to which the matrix is diagonal) are again calcu- lated from (109.5) and (87.1), this time with A = 0, S = 1. The non-zero elements are of the form (n.S)^ = VK7+i)/(2/+i)L ( n - s )/ +1 = VL//(2J+i)]. For a given /, the number K can take the values K = J, J±\. For the matrix elements V& , we find V J j = B e J(J+\)+* e , V£l = B e {J-\)J+* e {J+\)l{2J+\), V jtl = V i-l = a eV[/(/+l)]/(2/+l). We see that there are no transitions between states with K = J and those with K = J±.\. Hence one of the levels is simply E x = Vj. The other two (E 2 , E a ) are obtained by solving the quadratic secular equation formed from the matrix elements V J j~\, V J j%\, F/+i- Since we are here interested only in the relative position of the components of the triplet, we subtract the constant a e from all three energies E lt E 2 , E 3 . As a result we obtain *i =*./(/+!), E 2S = J B 6 (7 2 +/+l)-K±V[B e 2 (2/+l) 2 -a e J5 e +K 2 ]. In case b (a small), by considering three levels with the same K and different J (J = K, K±l), we again obtain the formulae of Problem 1. §85. Multiplet terms. Cases c and d Besides cases of a and b coupling and those intermediate between them, there are also other types of coupling. These originate as follows. The occurrence of the quantum number A is due ultimately to the electric interac- tion of the two atoms in the molecule, which results in the axial symmetry of the problem of determining the electron terms (this interaction in the molecule is called the coupling between the orbital angular momentum and the axis). The distances between terms with different values of A give a measure of the magnitude of this interaction. Previously we have tacitly supposed this interaction so strong that these distances are large both com- pared with the intervals in the multiplet splitting and compared with those in the rotational structure of the terms. There are, however, opposite cases where the interaction of the orbital angular momentum with the axis is com- parable with or even small compared with the other effects ; in such cases, of course, we cannot in any approximation speak of a conservation of the pro- jection of the orbital angular momentum on the axis, so that the number A is no longer meaningful. 308 The Diatomic Molecule §85 If the coupling of the orbital angular momentum with the axis is small in comparison with the spin-orbit coupling, we say that we have case c. It is found in molecules which contain an atom of a rare-earth element. These atoms are characterised by the presence of / electrons with uncompensated angular momenta ; their interaction with the axis of the molecule is weakened by the deep position of the / electrons in the atom. Cases intermediate be- tween the a and c types of coupling are found in molecules consisting of heavy atoms. If the coupling of the orbital angular momentum with the axis is small compared with the intervals in the rotational structure, we say that we have case d. This case is found for high rotational levels (with large J) in some elec- tron terms of the lightest molecules (H 2 , He 2 ). These terms are characterised by the presence in the molecule of a highly excited electron, whose interaction with the remaining electrons (or, as we say, with the "core" of the mole- cule) is so weak that its orbital angular momentum is not quantised along the axis of the molecule (whereas the "core" has a definite angular momen- tum Acore about the axis). As the distance r between the nuclei increases, the interaction between the atoms is diminished, and finally becomes small compared with the spin-orbit interaction within the atoms. Hence, if we consider the electron terms for fairly large r, we shall have case c. This must be borne in mind when ascertaining the relation between the electron terms of the molecule and the states of the atoms obtained as r -> oo. In §80 we have already discussed this relation, neglecting the spin-orbit interaction. When the fine structure of the terms is included, there arises also the question of the relation between the values J x and / 2 of the total angular momenta of the isolated atoms and the values of the quantum number Q, for the molecule. We shall give the results here, without reiterating arguments which are entirely similar to those of §80. If the molecule consists of different atoms, the possible values off |0| obtained on combining atoms with angular momenta J lt J 2 (J x ^ J 2 ) are given by the same table (80.1), in which we must put^i, J 2 m place of L ly L 2 , and jO| in place of A. The only difference is that, for half-integral Jx+J z , the smallest value of \C1 \ is not zero as shown in the table, but \. For integral 7i+/ 2 , on the other hand, there are 2/ 2 +l terms with Q. = 0, for which (as for S terms when the fine structure is neglected) we have to decide the question of sign. If J x and/ 2 are each half-integral, the number 2/ 2 +l is even, and there are equal numbers of terms, which we shall denote by + and 0~. If J x and/ 2 are both integral, however, then/ 2 +l terms are 0+ and J 2 are 0" (if (-lf^IJ, = 1) or vice versa (if {-\) J ^ J *IJ 2 = -1). If the molecule consists of similar atoms in different states, the resulting molecular states are the same as in the case of different atoms, the only f In adding the two total angular momenta J lt J t of the atoms to form the resultant angular momen- tum Q, the sign of Q is clearly immaterial. §86 Symmetry of molecular terms 309 difference being that the total number of terms is doubled, with each term appearing once as an even and once as an odd term. Finally, if the molecule consists of similar atoms in the same state (with angular momenta J x = J 2 = J), the total number of states is the same as in the case of different atoms, while their distribution in parity is such that, if / is integral and Q, is even, N g = N u +1; if J is integral and Q. is odd, N g = N u ; if J is half-integral and Q is even, N u = N g ; if J is half-integral and Q. is odd, N u = N g +1. All the 0+ terms are even and all the 0~ terms odd. As the nuclei approach, a coupling of type c usually passes into one of type a\. Here the following interesting circumstance may arise. As already mentioned, the term with A = belongs to case b, and as regards the classi- fication of case a this means that multiplet levels with different values of Q, (and the same A = 0) have the same energy; but such levels can occur on the approach of atoms which are in different fine-structure states. Thus it may happen that the same molecular term corresponds to different pairs of atomic fine-structure states. A similar situation may occur for terms with O = which, on the approach of the nuclei, become a molecular term with A # (and therefore S = —A). Such levels are doubly degenerate, since in case a the same energy corresponds to the terms 0+ and 0~ (which may arise from different pairs of atomic states). J §86. Symmetry of molecular terms In §78 we have already examined some symmetry properties of the terms of a diatomic molecule. These properties characterised the behaviour of the wave functions in transformations which leave the co-ordinates of the nuclei unaltered. Thus the symmetry of the molecule with respect to reflection in a plane passing through its axis brings about the difference between 2+ and S~ terms ; the symmetry with respect to a change in sign of the co-ordinates || of all the electrons (for molecules composed of like atoms) gives rise to the classification of terms into even and odd. These symmetry properties char- acterise the electron terms, and are the same for all rotational levels belonging to the same electron term. The states of the molecule, like those of any system of particles (see §30), are characterised by their behaviour with respect to inversion, i.e. a simul- taneous change in sign of the co-ordinates of all the electrons and the nuclei. For this reason, all the terms for the molecule can be divided into positive (whose wave functions are unaltered when the sign of the co-ordinates of the f The correspondence between the classification of terms of types a and c cannot be derived in a general form. Its derivation necessitates a consideration of the actual potential energy curves, taking into account the rule that levels of like symmetry cannot intersect (§79). t We here neglect what is called A-doubling (see §88). || The origin is supposed to be taken on the axis of the molecule, and half-way between the two nuclei. 310 The Diatomic Molecule §86 electrons and nuclei is reversed) and negative (whose wave functions change sign on inversion).f For A # 0, each term is doubly degenerate, on account of the two possible directions of the angular momentum about the axis of the molecule. As a result of inversion, the angular momentum itself does not change sign, but the direction of the axis of the molecule is reversed (since the atoms change places), and hence the direction of the angular momentum A relative to the axis of the molecule is reversed. Hence two wave functions belonging to the same energy level are transformed into each other, and from them we can always form a linear combination that is invariant with respect to inversion and one that changes sign under this transformation. Thus we obtain for each term two states, of which one is positive and the other negative. In practice, every term with A ^ is split, however (see §88), and so these two states correspond to different values of the energy. The S terms require special consideration to determine their sign. First of all, it is clear that the spin bears no relation to the sign of the term : the inversion operation changes only the co-ordinates of the particles, leaving the spin part of the wave function unaltered. Hence all the components of the multiplet structure of any given term have the same sign. In other words, the sign of the term depends only on K, and not on J.\ The wave function of the molecule is the product of the electron and nuclear wave functions. It has been shown in §82 that, in a S state, the motion of the nuclei is equivalent to that of a single particle, of orbital angular momentum K, in a centrally symmetric field U(r). Hence we can say that, when the sign of the co-ordinates is changed, the nuclear wave function is multiplied by (-1) K (see (30.7)). The electron wave function characterises the electron term, and to ascertain its behaviour under inversion we must consider it in a system of co-ordinates rigidly connected to the nuclei and rotating with them. Let x, y, z be a sys- tem of co-ordinates fixed in space, and £, rj, £ a. rotating system of co-ordinates in which the molecule is fixed. The direction of the axes of |, rj, £ is defined so that the £-axis coincides with the axis of the molecule from (say) nucleus 1 to nucleus 2, and the relative position of the positive directions of the axes of £, rj, £ is the same as in the system x, y, z (i.e. if the system x, y, z is left- handed, the system £, rj, £ is so too). As a result of the inversion operation, the direction of the axes of x, y, z is reversed, and the system changes from left-handed to right-handed. The system £, rj, £ must also become right- handed, but the £-axis, being rigidly connected to the nuclei, retains its former direction. Hence the direction of either one of the axes of £, rj must f We retain the customary terminology. It is unfortunate, however, since in the case of an atom the behaviour of the terms with respect to the operation of inversion is referred to as parity and not sign. The sign of which we are here speaking must not be confused with the + and — which are added as indices to S terms. J We recall that case b usually holds for S terms, and so it is necessary to use the quantum numbers K and /. §86 Symmetry of molecular terms 311 be reversed. Thus the operation of inversion in the fixed system of co- ordinates is equivalent in the moving system to a reflection in a plane passing through the axis of the molecule. Under such a reflection, however, the electron wave function of a 2+ term is unaltered, while that of a 2~" term changes sign. Thus the sign of the rotational components of a 2+ term is determined by the factor ( — 1)^; all the levels with even K are positive, while those with odd K are negative. For a 1r term, the sign of the rotational levels is deter- mined by the factor ( — 1) K+1 ; all levels with even K are negative, while those with odd K are positive. If the molecule consists of similar atoms, f its Hamiltonian is also invariant with respect to an interchange of the co-ordinates of the two nuclei. A term is said to be symmetric with respect to the nuclei if its wave function is un- altered when they are interchanged, and antisymmetric if its wave function changes sign. The symmetry with respect to the nuclei is closely related to the parity and sign of the term. An interchange of the co-ordinates of the nuclei is equivalent to a change in sign of the co-ordinates of all the particles (electrons and nuclei), followed by a change in sign of the co-ordinates of the electrons only. Hence it follows that, if the term is even and positive (or odd and negative), it is symmetric with respect to the nuclei. If, on the other hand, it is even and negative (or odd and positive), then it is antisymmetric with respect to the nuclei. At the end of §62 we have established a general theorem that the co-ordinate wave function of a system of two identical particles is symmetrical when the total spin of the system is even, and antisymmetrical when it is odd. If we apply this result to the two nuclei of a molecule composed of similar atoms, we find that the symmetry of a term is related to the parity of the total spin / obtained by adding the spins i of the two nuclei. The term is symmetric when i" is even, and antisymmetric when I is odd. % In particular, if the nuclei have no spin (i = 0), / is zero also ; hence the molecule has no antisymmetric terms. We see that the nuclear spin has an important indirect influence on the molecular terms, although its direct influence (the hyperfine structure of the terms) is quite unimportant. When the spin of the levels is taken into account, an additional degeneracy of the levels results. Again in §62, we have calculated the number of states with even and odd values of / that are obtained on adding two spins i. Thus, when i is half-integral, the number of states with even / is i(2«+l), and with odd / is (j+1)(2*"+1). From what was said above, we conclude that the ratio of the degrees g 8 , g a of the degeneracy || of symmetric and antisymmetric f The two atoms must be not only of the same element, but also of the same isotope. % Recalling the relation between the parity, sign and symmetry of terms, we conclude that, when the total spin I of the nuclei is even, the positive levels are even and the negative levels odd, and vice versa when J is odd. || The degree of degeneracy of a level is often referred to in this connection as its statistical weight. Formulae (86.1), (86.2) determine the ratio of the nuclear statistical weights of symmetric and anti- symmetric levels. 312 The Diatomic Molecule §87 terms for terms with half-integral i is gslga=il(i+V- (86.1) For integral /, we similarly find that this ratio is SJg a =(i+m (86.2) We have seen that the sign of the rotational components of a S + term is determined by the number (—1) K . Hence, for example, the rotational com- ponents of a 2+0 term for even K are positive, and therefore symmetric, while for odd K they are negative and consequently antisymmetric. Bearing in mind the results obtained above, we conclude that the nuclear statistical weights of the rotational components of a S + ff level with successive values of K take alternate values, in the ratios (86.1) or (86.2). A wholly similar situa- tion is found for E+ M , 1r g and S~ M levels. In particular, for i = the statistical weights of levels with even K for S + M and T,~ g terms, and of levels with odd K for T, + g and S~ M terms, are zero. In other words, in the electron states 2+ M , Tr g there are no rotational states with even K, and in 2+ , Hr u states there are none with odd K. Because of the extremely weak interaction of the nuclear spins with the electrons, the probability of a change in / is very small, even in collisions of molecules. Hence molecules differing in the parity of /, and accordingly having only symmetric or only antisymmetric terms, behave almost as differ- ent forms of matter. Such, for instance, are orthohydrogen and parahydro- gen; in the molecule of the former, the spins i = \ of the two nuclei are parallel (/ = 1), while in that of the latter they are antiparallel (/ = 0). §87. Matrix elements for the diatomic molecule In calculating the matrices of various quantities in the diatomic molecule, let us begin with the matrix elements for transitions between states with zero spin. Let n be a unit vector along the axis of the molecule. The vector n, re- garded as an operator, commutes with the operator of the energy of the electrons and with that of the vibrational energy, but not with the angular momentum K of the molecule. Hence the matrix of n is diagonal with respect to all the quantum numbers except K and M K (where by M K we denote the magnitude of the projection of the angular momentum K on a #-axis fixed in space). Omitting all suffixes but these two, we write the matrix elements in the form (K' 'M' 'k\o\KMk)- Their dependence on Mr is given directly by the general formulae (29.7), (29.9), in which L and M must be replaced by K and Mr, and the suffix n may be omitted. We denote the coefficients in these formulae by n^,, for the present case, so that, for example, (KM K \n z \KM K ) = «f . §87 Matrix elements for the diatomic molecule 313 To calculate the quantities n%, we start from the equations n . £ = A, n 2 = 1, written in the form (see (29.11)) \n+R.++\n-£.-+n z & z = A, n+n-+n z 2 = 1, and the commutation relation n g n+— n+rig = 0. Taking the diagonal matrix elements of these equations (the matrix elements of K being determined from the general formulae (27.12) with K, M K in place of L, M), we obtain, after some calculations which we here- omit, the following formulae for the quantities required (H. Honl and F. London 1925): H * = A Jw^) i W * _1 = (W * _l) * = iV[(K2 - A2)IK ^' (87 ' 1} For A = these formulae give n K = 0, n K ~i = WK; K K v » these correspond, as we should expect, to the matrix elements of a unit vector for motion in a centrally symmetric field (see §29, Problem 1). Next, let A be some vector physical quantity characterising the state of the molecule when the nuclei are fixed.f Let us first consider this quantity in the system of co-ordinates |, r\, £, which rotates with the molecule (the £-axis coinciding with the axis of the molecule). The results of §29 cannot here be applied in their entirety, since the angular momentum of the mole- cule with respect to the system of co-ordinates £ rj, £ (i.e. the electron angular momentum L) is not conserved ; only its ^-component A is conserved. The results concerning the selection rule for the quantum number A (M in §29) evidently remain fully valid, however. Thus the matrix elements of the vector A that are not zero are we denote by n the assembly of quantum numbers for the electron term, with the exception of A. If both the terms are S terms, we must also bear in mind the selection rule arising from the symmetry with respect to reflection in a plane passing through the axis of the molecule (the £-axis). Under such a reflection, the £-com- ponent of an ordinary (polar) vector is unchanged, while that of an axial vector changes sign. Hence we conclude that, for a polar vector, A c has non-zero matrix elements only for the transitions S+ -» S+ and S~ -> S~, while for t For example, the dipole moment or magnetic moment of the molecule. 314 The Diatomic Molecule §87 an axial vector the elements are non-zero for the transitions 2+ -> S~. We need not discuss the components A ^, A v , since for these no transitions with- out change of A are possible. If the molecule consists of similar atoms, there is also a selection rule re- garding parity. The components of a (polar) vector change sign under inver- sion. Hence their matrix elements are non-zero only for transitions between states of different parity (the reverse is true for an axial vector). In particular, all the diagonal matrix elements of the components of a polar vector vanish identically. The question arises how the matrix elements (87.2) are related to those of the same vector A in a fixed system of co-ordinates. In this system we can again use the general formulae (29.7), (29.9), which give the dependence of the matrix elements (nAKMK\A\n f A' K' M' k) on the quantum number Mr- The coefficients in these formulae are naturally denoted by A^j^, ; we have to relate these to the quantities (87.2). It is seen from (87.2) that there are matrix elements diagonal with respect to A only for the component along the axis of the molecule. Hence we can write the equation (nAKM K \A\nAK'M' K ) = {nAKM K \Ap\ri AK' M' K ). In particular, (nAKM K \A z \nAK'M K ) = (A^ A {KM K \n z \K'M K ). Separating out the dependence on M K , we hence have A nAK =n Kt A \*\ (87.3) where the n%, are determined by formulae (87.1). Thus we have found some of the relations which we desire. To find the remaining relations (for the components non-diagonal with respect to A), we notice that, since the quantity A refers to the molecule with fixed nuclei, the operators Ag, A v , A^ evidently commute with the vector n. The components of the vector A in the system x, y, z are linear combinations of the components in the system |, r], £, the coefficients in these combinations being functions of n x , n y , n z . Hence A x , A y , A z also commute with the vector n. In particular A z n z —n z 4 z = 0. Taking from this equation the matrix elements for the transitions A, K -> A— 1, K' with K' = K, K±\, we obtain three equations, from which the dependence on K of the required quantities A^^E lK , can be found. The coefficients in the resulting formulae can be related to the quantities (87.2) by comparing the matrix elements of the scalar A 2 , these being calcu- latedf in the two systems of co-ordinates x,y, z and £, rj, £. As a result, t The calculation is conveniently performed directly from the formula (29.12), where we must put A = B, replace L by K, and take n as w, A. §87 Matrix elements for the diatomic molecule 315 we obtain the following final formulae : A T.a-i.k ^m+^v^Vl^K+lXK+AXK-A+iyKiK+l)], n A 1**-i.k-x = ~*<^+^CSui V[(K+A)(K+A-1)IK], [(87.4) A 2tf-i!K = K^+iA^yiiK-AXK-A+iyK]. The components with A— 1 -* A are the complex conjugates of those given. Finally, we must find how the formulae we have obtained should be modi- fied for transitions between states with non-zero spin. Here it is important to know whether the states belong to case a or to case b. First, let the two states belong to case a. The unit vector n commutes with the spin vector S, and hence its matrix is diagonal with respect to the quantum numbers S and 2 (or, what is the same thing, S and Q., since £1 = A + 2, and this matrix is diagonal with respect to A also). The quantum numbers K and M K do not exist, and instead we have the total angular momentum / and its projection M on the #-axis. Instead of the relation n . K = A, which we used to derive (87.1), we now have n . J = Q.. Accord- ingly, we again obtain the same formulae (87.1), except that K and A must now be replaced by / and Q, respectively (we omit the diagonal suffix S). The same is true for any orbital vector A (i.e. one which does not depend on the spin). Such a vector commutes with S also, and hence its matrix is diagonal with respect to S and 2 ; if we use the quantum number Q. in place of 2, it changes together with A in the non-zero matrix elements (i.e. if A' = A± 1, then Q' = Cl± 1). The formulae (87.3) and (87.4) are unchanged except that we must add the suffixes Q. and Q', and everywhere (except the suffixes) replace K and A by / and CI. If the vector A depends on the spin, however, the selection rules are different. The vector S commutes with the orbital angular momentum, and also with the Hamiltonian, and hence its matrix is diagonal with respect to n and A; we omit these suffixes. It is, however, not diagonal with respect to 2 (or CI). The matrix elements of the components of S in the system $, rj, £ are deter- mined by formulae (27.13), with S and 2 in place of L and M, and then the transition to the system x, y, z is effected by the formulae (87.3), (87.4), where we must everywhere (including the suffixes) replace K and A by / and Q. Now let both states belong to case b. The calculation of the matrix ele- ments is here performed in two stages. First we consider the rotating mole- cule without taking into account the addition of the spin to the angular momentum K; the matrix elements are then determined by the same formulae (87.1)-(87.4). The vector A is supposed orbital, so that, like n, it commutes with S, and so the matrices are diagonal with respect to the quantum number S, which we omit from the suffixes. The angular momentum K is then 316 The Diatomic Molecule §88 added to S to form the total angular momentum J, and the transition to the new matrix elements is effected by the general formulae (109.3). The part of j x in these formulae is here taken by S, that of j 2 by K, and we write n, A in place of n±. If one of the states belongs to case a and the other to case b, the calculation of the matrix elements for transitions between the states is more involved ; we shall not here pause to consider this problem, f PROBLEMS Problem 1. Determine the Stark splitting of the terms for a diatomic molecule having a constant dipole moment, in the case where the term belongs to case a. Solution. The energy of a dipole d in an electric field <f is — d . <f . From considera- tions of symmetry, it is evident that the dipole moment of a diatomic molecule is directed along its axis ; d = dn, where d is a constant. Taking the direction of the field as the sr-axis, we obtain the perturbation operator in the form —dn z $. Determining the diagonal matrix elements of n z in accordance with the formulae derived above, we find that in case a the splitting of the levels is given by the formulaj AE Mj = -£dMjQIJ(J+l). Problem 2. The same as Problem 1, but for the case where the term belongs to case b (and A # 0). Solution. By the same method we have AE M = —SdMjK . 2K(K+1)J(J+1) Problem 3. The same as Problem 2, but for a *2 term. Solution. For A = the linear effect is absent, and we must go to the second approxi- mation of perturbation theory. In the summation in the general formula (38.9), it is sufficient to retain only those terms which correspond to transitions between rotational components of the electron term concerned; for other terms the energy differences in the denominators are large. Thus we find „*M KM k\< K - X > M K )\* UKMKlnzlK+^MK)]* AiiAf =d z z { 1 K \ Ek—Ek-i Ek—Ek+i where Ek = BK(K+1). A simple calculation gives « 2 K(K+1)-3M K * &E M = * B 2K(K+l)(2K-l)(2K+3) §88. A-doubling The double degeneracy of the terms with A ^ (§78) is in fact only approximate. It occurs only so long as we neglect the effect of the rotation t See E. Hill and J. van Vleck, Physical Review 32, 250, 1928. t It may seem that there is here a contradiction of the general assertion that there is no linear Stark effect (§76). In fact, of course, there is no contradiction, since the presence of a linear Stark effect is here due to the double degeneracy of the levels with CI ^ 0; the formula obtained is therefore applicable provided that the energy of the Stark splitting is large compared with that of what is called the A-doubling (§88). §88 A-doubling 317 of the molecule on the electron state (and also the higher approximations with respect to the spin-orbit interaction), as we have done throughout the above theory. When the interaction between the electron state and the rotation is taken into account, a term with A # is split into two levels close together. This phenomenon is called A-doubling (E. Hill and J. H. van Vleck, and R. de L. Kronig, 1928). To consider this effect quantitatively, we again begin with the singlet terms (S = 0). We have calculated (in §82) the energy of the rotational levels in the first approximation of perturbation theory, determining the diagonal matrix elements (i.e. the mean value) of the operator B(r)(R-L)«. To calculate the subsequent approximations, we must consider the elements of this operator that are not diagonal with respect to A. The operators & 2 and L 2 are diagonal with respect to A, so that we need consider only the oper- ator -2B&.L. The calculation of the matrix elements of R . L is conveniently effected by means of the general formula (29.12), in which we must put A = K, B = L ; the parts of L and M are taken by K and M K , while in place of n we must put n, A, where n denotes the assembly of quantum numbers (other than A) which determine the electron term. Since the matrix of the vector K, which is conserved, is diagonal with respect to «, A, while that of the vector L contains non-diagonal elements only for transitions in which A changes by unity (cf. what was said in §87 concerning an arbitrary vector A), we find, using formulae (87.4), (nAKMKlK.Lln^A-UKM^^^+iL^^VKK^AXK+l-A)^ (88.1) There are no non-zero matrix elements corresponding to any greater change in A. The perturbing effect of the matrix elements with A -> A— 1 can cause the appearance of an energy difference between states with ±A only in the 2Ath approximation of perturbation theory. Accordingly, the effect is pro- portional to B 2A , i.e. to (mjM) 2A , where M is the mass of the nuclei and m that of the electron. For A > 1, this quantity is so small that it is of no interest. Thus the A-doubling effect is of importance only for II terms (A = 1), which are considered below. For A = 1 we must go to the second approximation. The corrections to the eigenvalues of the energy can be determined from the general formula (38.9). In the denominators of the terms in the sum occurring in this equa- tion we have energy differences, of the form E nAK —E n , A _ 1K . In these differences, the terms containing K cancel, since, for a given distance r be- tween the nuclei, the rotational energy is the same quantity, B(r)K(K+l), for all the terms. Hence the dependence on K of the required splitting Ai? is entirely determined by the squared matrix elements in the numerators. 318 The Diatomic Molecule §88 Among these are the squared elements for transitions in which A changes from 1 to and from to — 1 ; these both give, by (88.1), the same depen- dence on K, and we find that the splitting of the 1 II term is of the form AE = constant x K(K+ 1), (88.2) where the constant is of the order of magnitude of B 2 ] e, e being the order of magnitude of the differences between neighbouring electron terms. Let us pass now to terms with non-zero spin ( 2 II and 3 II terms; higher values of S are not found in practice). If the term belongs to case b, the multiplet splitting has no effect on the A-doubling of the rotational levels, which is determined as before by formula (88.2). In case <z, however, the effect of the spin is important. Here each electron term is characterised by the number Q. as Well as A. If we simply replace A by —A, then Q = A+ 2 is changed, so that we obtain an entirely different term. The levels with A, Q. and —A, — Q are mutually degenerate. This degeneracy can here be removed not only by the effect, considered above, of the interaction between the orbital angular momentum and the rotation of the molecule, but also by the effect of the spin-orbit interaction. The con- servation of the projection Q. of the total angular momentum on the axis of the molecule is (if the nuclei are fixed) an exact conservation law, and so cannot be destroyed by the spin-orbit interaction; the latter can, however, change A and 2 (i.e. there are matrix elements for the corresponding transi- tions) in such a way that Q. remains unchanged. This effect, alone or in combination with the orbit-rotation interaction (which alters A but not 2), may cause A-doubling. Let us first consider the 2 II terms. For the 2 n i/2 term (A= 1, 2 =— \, Q, = |), the splitting is obtained on taking into account simultaneously the spin-orbit and orbit-rotation interactions, each in the first approximation. For the former gives the transition A = 1, 2= — ^ -> A = 0, 2= J, and then the latter converts the state A = 0, 2 = | into A = —1, 2 == \, which differs from the initial state by the signs of A and Q. being reversed. The mat- rix elements of the spin-orbit interaction are independent of the rotational quantum number /, while the dependence of those for the orbit-rotation interaction is determined by formula (88.1), in which (under the radical) we must replace K and A by / and Q. Thus we have for the A-doubling of a 2 II 1/2 term the expression AE 1/2 = constant x {J+Q, (88.3) where the constant is of the order of AB\ e. For a 2 n 3/2 term, on the other hand, the splitting can be found only in higher approximations, so that in practice &E 3/2 , = 0. Finally, let us consider 3 II terms. For a 3 II term (A = 1, 2 = —1), the splitting is obtained on taking into account the spin-orbit interaction in the second approximation (because of the transitions A = 1, 2 = —1 ->- A = 0, 2 = -> A = — 1, 2=1). Accordingly, the A-doubling in this case is §89 The interaction of atoms at large distances 319 entirely independent of /: &E = constant ~ A 2 je. (88.4) For a 3 I1 1 term, 2 = 0, and so the spin has no effect on the splitting; hence we again have a formula like (88.2), but with K replaced by / : AE X = constant x/(/+l). (88.5) For a 3 II 2 term, higher approximations are needed, so that we can suppose A£ 2 = 0. One of the levels of the doublet resulting from A-doubling is always posi- tive, and the other negative; we have already discussed this in §86. An investigation of the wave functions of the molecule enables us to establish the regularities of the alternation of positive and negative levels. Here we shall give only the results of the investigation.f It is found that if, for some value of J, the positive level is below the negative one, then in the doublet for /+ 1 the order is opposite, the positive level being above the negative one, and so on; the order varies alternately as the total angular momentum takes successive values. We are speaking here of case a terms; for case b, the same holds for successive values of the angular momentum K. PROBLEM Determine the A-splitting for a *A term. Solution. Here the effect appears in the fourth approximation of perturbation theory. Its dependence on K is determined by the products of the four matrix elements (88.1) for transitions with change of A : 2 -> 1, 1 -»- 0, -> —1, —1 -> —2. This gives AE = constant x(K-l)K(K+l)(K+ 2), where the constant is of order of -B 4 /* 3 - §89. The interaction of atoms at large distances Let us consider two atoms in S states which are at a great distance from each other (relative to their size), and determine the energy of their interaction. In other words, we shall discuss the determination of the form of the electron terms U n (r) when the distance between the nuclei is large. To solve this problem we apply perturbation theory, regarding the two isolated atoms as the unperturbed system, and the potential energy of their electrical interaction as the perturbation operator. As we know from electro- statics, the electrical interaction of two systems of charges at a large distance r apart can be expanded in powers of 1/r, and successive terms of this expansion correspond to the interaction of the total charges, dipole moments, quad- rupole moments, etc., of the two systems. For neutral atoms, the total charges are zero. The expansion here begins with the dipole-dipole interaction (~ 1/r 3 ); then follow the dipole-quadrupole terms (~ 1/r 4 ), the quadrupole- quadrupole (and dipole-octupole) terms (^ 1/r 5 ), and so on. f This may be found in E. Wigner and E. WiTMER, Zeitschrift fur Physik 51, 859, 1928. 320 The Diatomic Molecule §89 In the first approximation of perturbation theory, the required energy of the interaction of the atoms is determined as the diagonal matrix element of the perturbation operator, calculated with respect to the unperturbed wave functions of the system (expressed in terms of products of the unperturbed functions for the atoms).f In S states, however, the diagonal matrix elements, i.e. the mean values of the dipole, quadrupole, etc." moments, are zero ; this follows at once from considerations of symmetry, since the distribution of charges in an atom in the S state is spherically symmetrical on the average. Hence each of the terms of the expansion of the perturbation operator in powers of \\r gives zero in the first approximation of perturbation theory. J In the second approximation it is sufficient to restrict ourselves to the dipole interaction in the perturbation operator, since this decreases least rapidly as r increases, i.e. to the term V = [-d x . d 2 +3(d x . n)(d 2 . n)]/r\ (89.1) where n is a unit vector in the direction joining the two atoms. Since the non- diagonal matrix elements of the dipole moment are in general different from zero, we obtain in the second approximation of perturbation theory a non- vanishing result which, being quadratic in V, is proportional to 1/r 6 . The correction in the second approximation to the lowest eigenvalue is, as we know, always negative (§38). Hence we obtain for the interaction energy of atoms in their normal states an expression of the form || U(r) = -constant/r 6 , (89.2) where the constant is positive (F. London 1928). Thus two atoms in normal S states, at a great distance apart, attract each other with a force (— dU/dr) which is inversely proportional to the seventh power of the distance. The attractive forces between atoms at large distances are usually called van der Waals forces. These forces cause the appearance of minima on the potential energy curves of the electron terms even for atoms which do not form a stable molecule. These depressions, however, are very shallow (being only tenths or even hundredths of an electron-volt in depth) and lie at distances several times greater than the distances between atoms in stable molecules. If only one of the atoms is in the S state, the same result (89.2) is obtained for the interaction energy, since, for the first approximation to vanish, it is ■f Here we neglect the exchange effects, which decrease exponentially with distance (see §62, Problems). % This, of course, does not imply that the mean value of the interaction energy of the atoms is pre- cisely zero. It diminishes exponentially with distance, i.e. more rapidly than every finite power of 1/r, and hence each term of the expansion vanishes. This occurs because the expansion of the interaction operator in terms of the multipole moments involves the assumption that the charges of the two atoms are at a large distance r apart, whereas in quantum mechanics the electron density distribution has finite (though exponentially small) values even at large distances. || For brevity, we here and later omit the unimportant constant term in U(r), i.e. the value of U(co), which is the sum of the energies of the two isolated atoms. §89 The interaction of atoms at large distances 321 sufficient for the dipole (etc.) moment of only one atom to be zero. The constant in the numerator of (89.2) here depends, not only on the states of the two atoms, but also on their mutual orientation, i.e. on the value Q. of the projection of the angular momentum on the axis joining the atoms. If both atoms have non-zero orbital and total angular momenta, however, the situation is changed. The mean value of the dipole moment is zero in every state of the atom (§75). The mean values of the quadrupole moment in states with L # 0, / # or \ are not zero, however. Hence the quadrupole- quadrupole term in the perturbation operator gives a non-zero result even in the first approximation, and we find that the interaction energy of the atoms diminishes as the fifth, not the sixth, power of the distance : U(r) = constant/r 5 . (89.3) Here the constant may be either positive or negative, i.e. we may have either attraction or repulsion. As in the previous case, this constant depends not only on the states of the atoms, but also on the state of the system formed by the two atoms. A special case is the interaction of two similar atoms in different states. The unperturbed system (the two isolated atoms) has here an additional de- generacy due to the possibility of interchanging the states of the atoms. Accordingly, the correction in the first approximation will be given by the secular equation, in which the non-diagonal matrix elements of the perturba- tion appear as well as the diagonal ones. If the states of the two atoms have different parities, and angular momenta L differing by ± 1 or but not both zero (the same restriction being placed on J), then the non-diagonal matrix elements of the dipole moment for transitions between these states are in general not zero. Hence an effect in the first approximation is obtained from the dipole term in the perturbation operator. Thus the interaction energy of the atoms is here proportional to 1/r 3 : U(r) = constant/r 3 , (89.4) where the constant may have either sign. Usually, however, what is of interest is the interaction of the atoms aver- aged over all possible states of the system which they form (for given states of the atoms), including all possible orientations of the angular momenta of the atoms.-|- As a result of this averaging, all effects linear in the dipole or quadrupole moment of each atom (i.e. all effects in the first approximation of perturba- tion theory) vanish. The averaged interaction forces between atoms at large distances therefore always follow the law (89.2)J. t Such averaging is necessary, for example, in the problem of determining the interaction of atoms in a gas. t This law, derived on the basis of the non-relativistic theory, is Valid only so long as the retardation of electromagnetic interactions is unimportant. For this to be so, the distance r between the atoms must be small compared with c/cuon, where won are the frequencies of transitions between the ground state and the excited states of the atom. 322 The Diatomic Molecule §90 PROBLEM Derive a formula giving the van der Waals forces in terms of the matrix elements of dipole moments for two like atoms in S states. Solution. The answer is obtained by applying the general formula (38.9) of perturbation theory to the operator (89.1). On account of the isotropy of the atoms in the S state it is evident a priori that, on summation over all intermediate states, the squared matrix elements of the three components of each of the vectors di and d2 give equal contributions, while the terms which contain products of different components give zero. The result is T7/ , 6 ^-a (d z )on 2 (d z )on' 2 U(r) = > , r«^2E Q -E n -E n > n,n' where Eo and E n are the unperturbed values of the energies of the ground state and excited states of the atom. Since by hypothesis L = in the ground state, the matrix elements (rfz)on are non-zero only for transitions to P states (L = l).t Using formulae (29.7), we bring U(r) to the final form U(r)= > ^ relM n V , 3r6 ^ 2E -E nl -En'i n,n' where in the matrix suffixes nL the second suffix gives the value of L and the first represents the assembly of the remaining quantum numbers which determine the energy level.J §90. Pre-dissociation A basic premise of the theory of diatomic molecules as given in this chapter is the assumption that the wave function of the molecule falls into the product of an electron wave function (depending on the distance between the nuclei as a parameter) and a wave function for the motion of the nuclei. This sup- position amounts to neglecting, in the exact Hamiltonian of the molecule, certain small terms corresponding to the interaction of the nuclear and electron motions. When these terms are taken into account and perturbation theory is applied, transitions between different electron states appear. || Physically, the transi- tions between states of which at least one belongs to the continuous spectrum are of particular importance. Fig. 30 shows curves for the potential energy of two electron terms.ff The energy E' (the lower dashed line in Fig. 30) is the energy of some vibrational level of a stable molecule in the electron state 2. In state 1, this energy lies in the range of the continuous spectrum. In other words, in passing from state 2 to state 1 the molecule automatically disintegrates ; this phenomenon is called pre-dissociation.%% As a result of pre-dissociation, the state of the dis- crete spectrum corresponding to curve 2 has in reality a finite lifetime. This t The spin of the atoms is not involved here. % Calculation gives for the coefficient in U = — Ar~ 6 (with all quantities in atomic units) the values A = 6 - 5 for two hydrogen atoms and A = 1*6 for helium atoms. |( As well as the splitting of the levels by A-doubling (§88). tf Strictly speaking, these curves must represent the effective potential energy Uj in some given rotational states of the molecule. {J Curve 1 may have no minimum at all if it corresponds to purely repulsive forces between the atoms. §90 Pre-dissociation 323 means that the discrete energy level is broadened, i.e. acquires a certain width (see the end of §44). Fig If, on the other hand, the total energy E lies above the dissociation limit in both states (the upper dashed line in Fig. 30), the transition from one state to the other corresponds to what is called a collision of the second kind. Thus the transition 1 -> 2 signifies the collision of two atoms, as a result of which the atoms are left in excited states, and separate with diminished kinetic energy (for r -> oo, curve 1 passes below curve 2; the difference C/ 2 (oo) — U^oo) is the excitation energy of the atoms). Because of the large masses of the nuclei their motion is quasi-classical. The problem of determining the probability of the transitions under considera- tion is therefore of the kind discussed in §52. From the general considera- tions given there we can say that the transition probability will be mainly determined by the point at which the transition could occur classically. f Since the total energy of the system of two atoms (the molecule) is conserved in the transition, the condition for it to be "classically possible" is that the effective potential energies should be equal: Uji(r) = C/j 2 (r). On account of the conservation of the total angular momentum of the molecule also, the centrifugal energies are the same in the two states, and so this condition means that the potential energies are equal: Vl {r) = U 2 (r), (90.1) the angular momentum not being involved at all. If equation (90.1) has no real roots in the classically accessible region (where E > Uji, Uj2), the transition probability according to §52 is expo- nentially small.J Transitions occur with an appreciable probability only if t Or else by the point r — at which the potential energy becomes infinite. t A peculiar situation must occur in the case of a transition involving a molecular term which can arise from two different pairs of atomic states (see the end of §85), i.e. when the potential energy curve is, as it were, split into two branches with increasing distance. In this case the transition probability should be considerably greater, but the problem has not yet been discussed in the literature. 324 The Diatomic Molecule §90 the potential energy curves intersect in the classically accessible region (as shown in Fig. 30). Then the exponent in formula (52.1) is zero (and this formula is therefore, of course, invalid) ; accordingly, the transition probability is determined by a non- exponential expression which will be derived below. The condition (90.1) can then be interpreted as follows. If the potential (and total) energies are the same, so are the linear momenta. Hence the condition (90.1) may also be written in the form r i = r 2 , Pi = p*> (90.2) where p is the momentum of the relative radial motion of the nuclei, and the suffixes 1 and 2 refer to the two electron states. Thus we can say that the distance between the nuclei and their relative momentum remain unchanged at the instant when the transition occurs (this is called Franck and Condon's principle). Physically, this is due to the fact that the electron velocities are large compared with those of the nuclei, and "during an electron transition" the nuclei cannot noticeably change their position or velocity. It is not difficult to establish the selection rules for the transitions in ques- tion. First of all, there are two obvious exact rules. The total angular momentum / and the sign of the term (positive or negative; see §86) cannot change in a transition. This follows at once from the fact that the conserva- tion of the total angular momentum and of the behaviour of the wave function under inversion of the co-ordinate system are exact laws for any (closed) system of particles. Next, the rule which forbids (for molecules composed of similar atoms) transitions between states of unlike parity is very nearly accurate. For the parity of the state is uniquely determined by the nuclear spin and the sign of the term. The conservation of the sign of the term is an exact law, however, while the nuclear spin is very nearly conserved by virtue of the weakness of its interaction with the electrons. The requirement that there should be a point of intersection of the potential energy curves means that the terms must be of different symmetry (see §79). Let us consider transitions occurring in the first approximation of perturba- tion theory; the probability of transitions which occur only in higher approxi- mations is relatively small. First of all, we notice that the terms in the Hamiltonian which lead to the transitions in question are just those which cause the A-doubling of the levels. Among these terms are, firstly, terms representing the spin-orbit interaction. They are the product of two axial vectors, of which one is of spin character (i.e. is composed of the operators of the electron spins), and the other is of co-ordinate character ; we emphasise, however, that these vectors are not simply the vectors S and L. Hence they have non-zero matrix elements for transitions in which S and A change by 0, ±1. The case where AS and AA are both zero (and A # 0) must be omitted, since the symmetry of the term would then be unchanged in the transition. The transition between two 2 terms is possible if one of them is §90 Pre-dissociation 325 a 2+ term and the other a S~ term; an axial vector has non-zero matrix elements only for transitions between 2+ and 2~ (see §87). The term in the Hamiltonian which corresponds to the interaction between the rotation of the molecule and its orbital angular momentum is proportional to j . L. Its matrix elements are non-zero for transitions with A A = ±1 without change of spin (only the ^-component of the vector, i.e. L v has ele- ments with AA = 0, but L c is diagonal with respect to the electron states). As well as the terms we have considered, there is also a perturbation due to the fact that the operator of the kinetic energy of the nuclei (i.e. the operator of differentiation with respect to the co-ordinates of the nuclei) acts, not only on the wave function of the nuclei, but also on the electron function, which depends on r as a parameter. The corresponding terms in the Hamiltonian are of the same symmetry as the unperturbed Hamiltonian. Hence they can lead only to transitions between electron terms of like symmetry, the prob- ability of which is negligible in view of the non-intersection of these terms. Let us go on to the actual calculation of the transition probability. For definiteness, we shall consider a collision of the second kind. According to the general formula (43.1), the required probability is given by the expression 2 w h = -T /xnuc,2*^)Xnuc,l dy (90.3) where x = f'/'nuc ('/w Denl g the wave function of the radial motion of the nuclei) and V(r) is the perturbing energy; we have taken, as the quantity v in (43.1), the energy E and integrated with respect to it. The final wave func- tion Xnuc,2 must be normalised by the delta function of energy. The quasi- classical function (47.4a), thus normalised, is *"WiHi/* d '-4 (904) a, The normalising factor is determined by the rule given at the end of §21. The wave function of the initial state can be written in the form ^-^""fi/**-*-}- (90 - 5) It is normalised so that the current density is unity in each of the two travelling waves into which the stationary wave (90.5) can be resolved; v x and v 2 are the velocities of the relative radial motion of the nuclei. On substituting these functions in (90.3), we obtain the dimensionless transition probability w. It can be regarded as the transition probability for the nuclei to pass twice the point r = r (the point of intersection of the levels). It must be borne in mind that the wave function (90.5) corresponds, in a certain sense, 326 The Diatomic Molecule §90 to a double passage through this point, since it contains both the incident and the reflected travelling waves. The matrix element of V(r), calculated with respect to the functions (90.4), (90.5), contains in the integrand a product of cosines, which can be written in terms of the cosines of the sum and difference of the arguments. On integrating near the point r = r where the terms intersect, only the second cosine is important, so that V(r) dr 4 1 f T 1 T , 1 C lV(r)dr = id cos t Px dr p 2 dr — The integral rapidly converges as we move away from the point of intersec- tion. Hence we can expand the argument of the cosine in powers of £ = r — r and integrate over £ from — oo to + oo (replacing the slowly varying coeffi- cient of the cosine by its value at r = r ). Bearing in mind that, at the point of intersection, p x = p 2 , we find r r where S is the value of the difference of the integrals at the point r = r . The derivative of the momentum can be expressed in terms of the force P= -dU/dr: differentiating the equation p^jlfi + U x = p^jljn + U 2 (where H is the reduced mass of the nuclei), we have v x dpjdr—v^ dpjdr = ^—F^. Thus r r Jftdr-J Jf -pi p % dr~S 9 +-±- -e, 2v where v is the common value of v x and v 2 at the point of intersection. The integration is effected by means of the well-known formula oo f cos(a+j8F) d£ = /^cos( a +^r), and as a result we have 8nV 2 /S, w = cos hv\F 2 -F x \ <r*> (90.6) The quantity S Q jh is large and varies rapidly with the energy E. Hence, on averaging over even a small interval of energy, the squared cosine can be replaced by its mean value. As a result we obtain the formula to = 4t7F 2 /^|F 2 -F 1 | (90.7) §90 Pre-dissociation 327 (L. Landau 1932). All the quantities on the right-hand side of the equation are taken at the point of intersection of the potential-energy curves. In the application to pre-dissociation, we are interested in the probability of the disintegration of the molecule in unit time. In this time, the nuclei in their vibrations pass 2(o>/2tt) times through the point r = r (where co is the angular frequency of the vibrations). Hence the required pre-dissociation probability is obtained by multiplying w (the probability for a double passage) by o>/2tt, i.e. it is IV^lhv^-F^. (90.8) The following remark must be made concerning these calculations. In speaking of the intersection of terms, we have had in mind the eigenvalues of the "unperturbed" Hamiltonian i? of the electron motion in the molecule ; in this, the terms V which lead to the transitions concerned are not taken into account. If we include these terms in the Hamiltonian, the intersection of the terms becomes impossible, and the curves move apart slightly, as shown in Fig. 3 1 . This follows from the results of §79 when regarded from a slightly different point of view. Let U JX (r) and U J2 (r) be two eigenvalues of the operator 3 (in which r is regarded as a parameter). In the region near the point r where the curves U JX {r) and Uj z (r) intersect, to determine the eigenvalues U(r) of the perturbed operator i? +^ we must use the metn °d given in §79, as a result of which we obtain the formula u(r) = K^x+^+^n+^i V[K^i-^2+^ii-^2 2 ) 2 +l^i 2 | 2 ]; (90.9) the matrix elements V Uf F" 22 , F 12 , like Uj-^ and U J2 , are functions of r. The interval between the two levels is now At/ = V[(^/l-^/2+^ll-^22) 2 + 4ir i2 |2]. (90.10) 328 The Diatomic Molecule §90 Hence it is clear that, if there are transitions between the two states (i.e. the matrix element V X2 is not zero), the intersection of the levels disappears. The least distance between the curves is now A=2|F 12 |. The formulae obtained above for the transition probability are applicable only so long as the "divergence" of the curves is fairly small. If the latter becomes considerable, the transition probability cannot be calculated by ordinary perturbation theory. To examine this problem, we use the following method (C. Zener 1932). Let i/»i, «/r 2 be the wave functions of the electron states corresponding to the "unperturbed" terms Ujx and U J2 , i.e. these functions are solutions of the equations #o^i = Ujifa, fi ip 2 = Uj 2 xft 2 . Let us seek the solution of the perturbed wave equation iH&P/dt =(# +P)Y in the form T = b x {t)^ x +b 2 {t)^ 2 . (90.11) Substituting this expression in the wave equation, multiplying the latter firstly by ip x and secondly by «/r 2 , and integrating, we obtain two equations for the functions b x {t) and b 2 (t): ihdbjdt = U Jx b x +Vb 2 , ifrdbjdt = U j2 b 2 +Vb x ; (90.12) we here include V xx , V 22 in U JX , U J2 , and denote V X2 by V(r) simply, in agree- ment with the notation in the preceding formulae. We consider the motion of the nuclei quasi-classically. Accordingly, the variables r and * are related by drjdt = v, where v is the classical velocity of the nuclei. Near the point r where the curves of Uj^r) and U J2 (r) intersect (shown by the dashed line in Fig. 31), we can expand Uj x and U J2 as series of powers of £ — r—r , writing V Jx = Uj-F Jx £, U J2 = Uj-Fj^ (90.13) where Uj is the common value of U JX and U J2 at the point r = r , and we have introduced the notation Fj = -{dUjldr) u . Introducing also new un- knowns a x , a % in place of b x , b 2 by means of b x = ajerVmvj, b 2 = a 2 e-QW u J f , (90.14) and replacing the differentiation with respect to t by one with respect to $ (d/dt = t;d/d£), we obtain from equations (90.12) ihadajdg = -F Jx £ ai +Va 2 , ihv da 2 /d$ = - Fj 2 fr 2 + Va lt (90.15) §90 Pre-dissociation 329 where v and V may be taken with sufficient accuracy to have their values at the point of intersection. If we solve equations (90.15) with the boundary condition a x = 1, a % — as £j -> +oo, then |0i(-oo)| 2 determines the probability that, as the nuclei pass through the point £ = 0, the molecule remains in the electron state lf indicating a transition from the curve 12' to the curve 21' (see Fig. 31). Similarly, \a^ - oo) | 2 = 1 - \a x { - oo) | 2 is the probability of a transition to the electron state «^ 2 , i.e. the probability that the molecule remains on the curve 12'. The transition from curve 1 to curve 2 (as £ -> + oo) in a double passage through the point of intersection can be effected in two ways: either by 1 -+ V -> 2 (as the nuclei approach, the transition from curve 12' to curve 21' occurs, and as they recede the molecule remains on the curve 21'), or by 1 _> 2' -> 2. Hence the required probability for such a transition is W =2K(-oo)|*{lH*i(-°°)l 2 }- We shall not pause here to explain the manner of solution of equations (90.15) (they reduce to a single equation of the second order, which can be solved by Laplace's method), but give only the final result :f (the difference Fj^—F^ is replaced by the equal difference F 2 —F x ). Thus w = 2e- 2vV ^ F t- F ^{l—e- 2nV ^ nvlF *- F ^}. (90.16) We see that the probability of the transition in question is small in two limiting cases, when V is fairly small and fairly large. For V 2 <^ hv\F 2 —F x \, formula (90.16) becomes (90.7). Finally, let us consider the phenomenon, akin to pre-dissociation, of what are called perturbations in the spectra of diatomic molecules. If two discrete molecular levels E x and E 2 corresponding to two intersecting electron terms are close together, the possibility of a transition between the two electron states results in a displacement of the levels. According to the general formula (79.4) of perturbation theory, we have for the displaced levels the expression K£i+£ 2 )± V[m-E z y+\ v 12tRUC \*], (90.17) where Fi 2fnuc is the matrix element of the perturbation for the transi- tion between the molecular states 1 and 2; the matrix elements V llnuc and Voo must, of course, be included in & and E 2 . From this formula we see that the two levels are moved apart, being displaced in opposite directions (the higher level is raised and the other lowered). The amount of the dis- placement is the greater, the smaller the difference \E X —E 2 \. The matrix element r i2nuc is calculated in exactly the same way as for determining the probability of a collision of the second kind. The only t See C. Zener {Proceedings of the Royal Society A 137, 696, 1932). 330 The Diatomic Molecule §90 difference is that the wave functions Xnuc§1 and Xnuc2 belong to the discrete spectrum, and hence must be normalised to unity. ' According to (48.3) we have for these functions w ~y;? "fiiJ* dr -H A comparison with formulae (90.3) to (90.5) shows that the matrix element Pi2,nuc ner e considered is related to the transition probability w for a double passage through the point of intersection by I ^i2,nud 2 = w(^<o 1 /27r)(^co 2 /27r). (90.18) PROBLEMS Problem 1. Determine the total effective cross-section for collisions of the second kind, as a function of the kinetic energy E of the colliding atoms, for transitions pertaining to the spin-orbit interaction. Solution. On account of the quasi-classical motion of the nuclei, we can introduce the concept of the impact parameter p (the distance at which the nuclei would pass if there were no interaction between them) and define the effective cross-section da as the product of the "target area" 2np dp and the transition probability w(p) per collision, f The total effective cross-section a is obtained by integrating with respect to p. For spin-orbit interaction, the matrix element V(r) is independent of the angular momen- tum M of the colliding atoms. We write the velocity v at the point r = r , where the curves intersect, in the form v = ViWME-U-MiptirS)-] = VK2/ix)(E-U-fElr *)]. Here U is the common value of U x and U 2 at the point of intersection, p. is the reduced mass of the atoms, and the angular momentum M = ppvao, where v«> is the relative velocity of the atoms at infinity. The zero of energy is chosen so that the interaction energy of the atoms in the initial state is zero at infinity; then E = ip,v m *. Substituting this expression in (90.7), we find 8tt 2 F 2 pdp da = 2vp dp . w = *l*r-*il V[2(E-U-p*E/r *)lp.] ust be taken from zero uj ave The integration with respect to p must be taken from zero up to the value for which the velocity v vanishes. As a result we have HlFt-FJl E Problem 2. The same as Problem 1, but for transitions pertaining to the interaction be- tween the rotation of the molecule and its orbital angular momentum. t Cf. Mechanics, §18. §90 Pre-dissociation 331 Solution. The matrix element V is of the form V(r) = MD\[ir % , where D(r) is the matrix element of the electron orbital angular momentum. By the same method as in Problem 1 we obtain 16V27r 2 Z> 2 (E-U)*'* 3hVn\F*-Fi\ E ' Problem 3. Determine the transition probability for energies E close to the value Uj of the potential energy at the point of intersection. Solution. For small values of E-Uj, formula (90.7) is inapplicable, since the velocity v of the nuclei cannot be regarded as constant near the point of intersection, and hence it cannot be taken outside the integral as it was in deriving (90.7). Near the point of intersection we replace the curves of Uj u Uj t by the straight lines (90.13). The wave functions Xnuc,i and Xn™,2 in this region are wave functions of one-dimensional motion in a homogeneous field (§24). The calculations are conveniently effected by means of wave functions in the momentum representation. The wave function normalised by the delta function of energy is of the form (see §24, Problem) a 9 = V{2*h\F Jz \) exp \-^ri( E - u j)P~P z l 6 ri } • while the wave function normalised to unit current density in the incident and reflected waves is obtained by multiplying by \/(2irH) : 1 ( i ■* --^n ex ^\jw- [{E - Uj)p - pzl6tl] V\ F Ji\ WJi On integrating, the perturbing energy (matrix element) V may again be taken outside the integral, replacing it by its value at the point of intersection; °° 2tt\ r 2 to = — \V c^a^dp h • J — 00 As a result we obtain 47rF 2 (2/Li) 2 / 3 ^3 r /2/A 1/3 / 1 1 \ 2/3 "l where $(£) is the Airy function (see §b of the Mathematical Appendices). For large E—Uj, this formula reduces to (90.7). CHAPTER XII THE THEORY OF SYMMETRY §91. Symmetry transformations The classification of terms in the polyatomic molecule is fundamentally related to its symmetry, as in the diatomic molecule. Hence we shall begin by examining the types of symmetry which a molecule can have. The symmetry of a body is determined by the assembly of all those re- arrangements after which the body is unaltered; these rearrangements are called symmetry transformations. Any possible symmetry transformation can be represented as a combination of one or more of the three fundamental types of transformation. These three essentially different types are: the rotation of the body through a definite angle about some axis, the reflection of it in some plane, and the parallel displacement of the body over some distance. Of these, the last evidently is applicable only to an infinite medium (a crystal lattice). A body of finite dimensions (in particular, a molecule) can be symmetrical only with respect to rotations and reflections. If the body is unaltered on rotation through an angle lirjn about some axis, then that axis is said to be an axis of symmetry of the nth order. The number n can take any integral value: n = 2, 3, ... . The value n = 1 corresponds to a rotation through an angle of 2n or, what is the same thing, of 0, i.e. it corresponds to an identical transformation. We shall symbolically denote by C n the operation of rotation through an angle 2rt\n about a given axis. Repeating this operation two, three, ... times, we obtain rotations through angles 2(2?r/«), 3(27r/«), ... , which also leave the body unaltered; these rotations may be denoted by C n 2 , C n 3 , .... It is obvious that, if p divides n, C n p = C n/P . (91.1) In particular, performing the rotation n times, we return to the initial position, i.e. we effect an identical transformation. The latter is customarily denoted by E, so that we can write C n n = E. (91.2) If the body is left unaltered by a reflection in some plane, this plane is said to be a plane of symmetry. We shall denote by the symbol a the operation of reflection in a plane. It is evident that a double reflection in the same plane is the identical transformation : <r 2 = E. (91.3) 332 §91 Symmetry transformations 333 A simultaneous application of the two transformations (rotation and reflection) gives what are called the rotary-reflection axes. A body has a rotary-reflection axis of the nth order if it is left unaltered by a rotation through an angle lirjn about this axis, followed by a reflection in a plane perpendicular to the axis (Fig. 32). It is easy to see that this is a new form 1 oP Kl Fig. 32 of symmetry only when n is even. For, if n is odd, an w-fold repetition of the rotary-reflection transformation would be equivalent to a simple reflection in a plane perpendicular to the axis (since the angle of rotation is 2tt, while an odd number of reflections in the same plane amounts to a simple reflection). Repeating this transformation a further n times, we have as a result that the rotary-reflection axis reduces to the simultaneous presence of an axis of symmetry of the nth order and an independent plane of symmetry perpen- dicular to this axis. If, however, n is even, an w-fold repetition of the rotary- reflection transformation returns the body to its initial position. We denote the rotary-reflection transformation by the symbol S n . Denoting by a h a reflection in a plane perpendicular to a given axis, we can put, by definition, S n = C n c h = a h C n \ (91.4) the order in which the operations C n and a h are performed clearly does not affect the result. An important particular case is a rotary-reflection axis of the second order. It is easy to see that a rotation through an angle tt, followed by a reflection in a plane perpendicular to the axis of rotation, is the inversion transformation, whereby a point P of the body is carried into another point P', lying on the continuation of the line which joins P to the intersection O of the axis and the plane, and such that the distances OP and OP' are the same. A body symmetrical with respect to this transformation is said to have a centre of symmetry. We shall denote the operation of inversion by /, so that we have I=S 2 = C 2 c h . (91.5) It is also evident that Ia h = C 2 , 7C 2 = a h \ in other words, an axis of the 334 The Theory of Symmetry §91 second order, a plane of symmetry perpendicular to it and a centre of sym- metry at their point of intersection are mutually dependent: if any two of these elements are present, the third is automatically present also. We shall now give various purely geometrical properties of rotations and reflections which it is useful to bear in mind in studying the symmetry of bodies. A product of two rotations about axes intersecting at some point is a rotation about some third axis also passing through that point. A product of two reflections in intersecting planes is equivalent to a rotation ; the axis of this rotation is evidently the line of intersection of the planes, while the angle of rotation is easily seen, by a simple geometrical construction, to be twice the angle between the two planes. If we denote a rotation through an angle about an axis by C(<f), and reflections in two planes passing through that axis by the symbols! a v and a' v , the above statement can be written as a v a' v = C(2cf>), (91.6) where <j> is the angle between the two planes. It must be noted that the order in which the two reflections are performed is not immaterial. The trans- formation a v o' v gives a rotation in the direction from the plane of a' v to that of a v ; on interchanging the factors we have a rotation in the oppo- site direction. Multiplying equation (91.6) on the left by a v , we obtain a' v = G v C(2<f>); (91.7) in other words, the operation of rotation, followed by reflection in a plane passing through the axis, is equivalent to a reflection in another plane intersecting the first at half the angle of rotation. In particular, it follows from this that an axis of symmetry of the second order and two mutually perpendicular planes of symmetry passing through it are mutually dependent ; if two of them are present, so is the third. We shall show that the product of rotations through an angle tt about two axes intersecting at an angle <f> (Oa and Ob in Fig. 33) is a rotation through p' Fig. 33 "f The suffix v customarily denotes a reflection in a plane passing through a given axis (a "vertical' plane), and the suffix h a reflection in a plane perpendicular to the axis (a "horizontal" plane). §92 Transformation groups 335 an angle 2^ about an axis perpendicular to the first two (PP' in Fig. 33). For it is obvious that the resulting transformation is also a rotation; after the first rotation (about Oa) the point P is carried into P', and after the second (about Ob) it returns to its original position. This means that the line PP r remains fixed, and is therefore an axis of rotation. To determine the angle of rotation, it is sufficient to note that, in the first rotation, the axis Oa remains fixed, while after the second it takes the position Oa\ which makes an angle 24 with Oa. In the same way we can see that, when the order of the two transformations is reversed, we obtain a rotation in the opposite direction. Although the result of two successive transformations in general depends on the order in which they are performed, in some cases the order of opera- tions is immaterial : the transformations commute. This is so for the following transformations : (1) Two rotations about the same axis. (2) Two reflections in mutually perpendicular planes (equivalent to a rotation through it about their line of intersection). (3) Two rotations through rr about mutually perpendicular axes (equivalent to a rotation through -it about the third perpendicular axis). (4) A rotation and a reflection in a plane perpendicular to the axis of rotation. (5) Any rotation or reflection and an inversion with respect to a point lying on the axis of rotation or in the plane of reflection; this follows from (1) and (4). §92. Transformation groups The set of all the symmetry transformations for a given body is called its symmetry transformation group (or simply its symmetry group). Hitherto we have spoken of these transformations as geometrical rearrangements of the body. However, in quantum-mechanical applications it is more convenient to regard symmetry transformations as transformations of the co-ordinates which leave the Hamiltonian of the system in question invariant. It is obvious that, if the system is left unaltered by some rotation or reflection, the cor- responding transformation does not change its Schrodinger's equation. Thus we shall speak of a transformation group with respect to which a given Schrodinger's equation is invariant.! f This point of view enables us to include in our considerations not only the rotation and reflection groups discussed here, but also other types of transformation which leave Schrodinger's equation unaltered. These include the interchange of the co-ordinates of identical particles forming part of the system considered (a molecule or atom). The set of all possible permutations of identical particles in a given system is called its permutation group (we have already met these permutations in §63). The general properties of groups given below apply to permutation groups also; we shall not pause to study this type of group in more detail here. The following remark should be made concerning the notation which we use in this chapter. Sym- metry transformations are essentially operators just like those which we consider all through the book (in particular, we have already considered the inversion operator in §30). They ought, therefore, to be denoted by letters with circumflexes. We do not do this, in view of the generally accepted notation, and because this omission cannot lead to misunderstandings in the present chapter. For the same reason we denote the identical transformation by the customary symbol E, and not by 1, which would correspond to the notation in the other chapters. 336 The Theory of Symmetry §92 Symmetry groups are conveniently studied with the help of the general mathematical techniques of what is called group theory, the fundamentals of which we shall explain below. At first we shall consider groups, each of which contains a finite number of transformations (known as finite groups). Each of the transformations forming a group is said to be an element of the group. Symmetry groups have the following important properties. Each group contains the identical transformation E (called the unit element of the group). The elements of a group can be "multiplied" by one another; by the product of two (or more) transformations we mean the result of applying them in succession. It is obvious that the product of any two elements of a group is also an element of that group. For the multiplication of elements we have the associative law (AB)C — A(BC), where A, B, C are elements of a group. There is evidently no general commutative law; in general, AB ^ BA. For each element A of a group there is in the same group an inverse element A -1 (the inverse transformation), such that AA~ X = E. In some cases an element may be its own inverse ; in particular, E _1 = E. It is evident that mutually inverse elements A and A -1 commute. The element inverse to the product AB of two elements is (AB)- 1 = Br*-A~\ and similarly for the product of a greater number of elements ; this is easily seen by effecting the multiplication and using the associative law. If all the elements of a group commute, the group is said to be Abelian. A particular case of Abelian groups is formed by what are called cyclic groups. By a cyclic group we mean a group, all of whose elements can be obtained by raising one of them to successive powers, i.e. a group consisting of the elements A, A\ A\ ... , A n = E, where n is some integer. Let G be some group.f If we can separate from it some set of elements H such that the latter is itself a group, then the group H is called a sub-group of the group G. A given element of a group may appear in several of its sub-groups. By taking any element A of a group and raising it to successive powers, we finally obtain the unit element (since the total number of elements in the group is finite). If n is the smallest number for which A n = E, then n is called the order of the element A, and the set of elements A, A 2 , ... , A n = E is called the period of A. The period is denoted by {A} ; it is itself a group, i.e. it is a sub-group of the original group, and is cyclic. In order to find whether a given set of elements of a group is a sub-group of it, it is sufficient to find whether, on multiplying any two of its elements, we obtain another element of the set. For in that case we have, together with t We shall denote groups by bold italic letters. §92 Transformation groups 337 each element A, all its powers, including A n ~ x (where n is the order of A), which is the inverse of A (since A n ~ x A = A n = E); and there will obviously be a unit element. The total number of elements in a group is called its order. It is easy to see that the order of a sub-group is a factor of the order of the whole group. To show this, let us consider a sub-group H of a group G, and let G x be some element of G which does not belong to H. Multiplying all the elements of H (on the right, say) by G ly we obtain a set (or complex, as it is called) of elements, denoted by HG X . All the elements of this complex clearly belong to the group G. However, none of them belongs to H; for, if for any two elements H a , H b belonging to H we had H a G x = H b , it would follow that G x = H a ~ x H b , i.e. G x would also belong to the sub-group H, which is contrary to hypothesis. Similarly we can show that, if G 2 is an element of G not belonging to H or to HG X , none of the elements of the complex HG 2 will belong to H or to HG V Continuing this process, we finally exhaust all the elements contained in the finite group G. Thus all the elements are divided among the complexes H, HG lt HG 2 , ... , HG m (where m is some integer), each of which contains h elements, h being the order of the sub-group H. Hence it follows that the order g of the group G is g = km, and this proves the theorem. If the order of a group is a prime number, it follows at once from the above that the group has no sub-groups (except itself and E). The converse theorem is also valid: a group having no sub-groups is of prime order and in addition must be cyclic (since otherwise it would contain elements whose period would form a sub-group). We shall now introduce the important concept of conjugate elements. Two elements A and B are said to be conjugate if A = CBC- 1 , where C is also an element of the group ; multiplying this equation on the right by C and on the left by C _1 , we have the converse equation B = C~ X AC. An important property of conjugate elements is that, if A is conjugate to B, and B to C, then A is conjugate to C; for, if B = P- X AP, C = Q~ X BQ (P and Q being elements of the group), it follows that C = (P0~M(P0. For this reason we can speak of sets of conjugate elements of a group. Such sets are called classes of the group. Each class is completely determined by any one element A of it ; for, given A, we obtain the whole class by forming the products GAG~ X , where G is successively every element of the group (of course, this may give each element of the class several times). Thus we can divide the whole group into classes ; each element of the group can clearly appear in only one class. The unit element of the group is a class by itself, since for every element of the group GEG~ X = E. If the group is Abelian, each of its elements is a class by itself; since all the elements, by definition, 338 The Theory of Symmetry §93 commute, each element is conjugate only to itself. We emphasise that a class of a group (not being E) is not a sub-group of it; this is evident from the fact that it does not contain a unit element. All the elements of a given class are of the same order. For, if n is the order of the element A (so that A n = E), then for a conjugate element B = CMO 1 we have {CAC- X ) n = CA»C~ X =* E. Let H be a sub-group of G, and G x an element of G not belonging to H. It is easy to see that the set of elements GiHG-f 1 has all the properties of a group, i.e. it also is a sub-group of the group G. The sub-groups H and G x HG-f x are said to be conjugate; each element of one is conjugate to one element of the other. By giving G t various values, we obtain a series of conjugate sub-groups, which may partly coincide. It may happen that all the sub-groups conjugate to H are H itself. In this case H is called a normal divisor of the group G. Thus, for example, every sub-group of an Abelian group is clearly a normal divisor of it. Let us consider a group A with n elements A, A', A", ... , and a group B with m elements B, B', B", ... , and suppose that all the elements of A (apart from the unit E) are different from those of B but commute with them. If we multiply every element of group A by every element of group B, we obtain a set of nm elements, which also form a group. For, for any two elements of this set we have AB . A'B' = AA' . BB' = A"B", i.e. another element of the set. The group of order nm thus obtained is denoted by A xB, and is called the direct product of the groups A and B. Finally, we shall introduce the concept of the isomorphism of groups. Two groups A and B of the same order are said to be isomorphous if we can establish a one-to-one correspondence between their elements, such that, if the element B corresponds to the element A, and B' to A', then B" = BB' corresponds to A" = AA' . Two such groups, considered in the abstract, clearly have identical properties, though the actual meaning of their elements may be different. §93. Point groups Transformations which appear in the symmetry group of a body of finite dimensions (in particular, a molecule) must be such that at least one point of the body remains fixed when any of these transformations is applied. In other words, all axes and planes of symmetry of a molecule must have at least one common point of intersection. For a successive rotation of the body about two non-intersecting axes or a reflection in two non-intersecting planes results in a translation of the body, which obviously cannot then be left unaltered. Symmetry groups having the above property are called point groups. Before going on to construct the possible types of point group, we shall explain a simple geometrical procedure whereby the elements of a group may be easily divided into classes. Let Oa be some axis, and let the element A of the group be a rotation through a definite angle about this axis. Next, let G §93 Point groups 339 be a transformation (rotation or reflection) in the same group, which on being applied to the same axis Oa carries it to the position Ob. We shall show that the element B = GAG' 1 then corresponds to a rotation about the axis Ob through the same angle as that of the rotation about Oa to which the element A corresponds. For, let us consider the effect of the transformation GAG' 1 on the axis Ob itself. The transformation G _1 inverse to G carries the axis Ob to the position Oa, so that the subsequent rotation A leaves it in this position; finally, G carries it back to its initial position. Thus the axis Ob remains fixed, so that B is a rotation about this axis. Since A and B belong to the same class, their orders are the same; this means that they effect rotations through the same angle. Thus we reach the result that two rotations through the same angle belong to the same class if there is, among the elements of the group, a transformation whereby one axis of rotation can be carried into the other. In exactly the same way, we can show that two reflections in different planes belong to the same class if some transformation in the group carries one plane into the other. The axes or planes of symmetry whose directions can be carried into each other are said to be equivalent. Some additional comments are necessary in the case where both rotations are about the same axis. The element inverse to the rotation C n k (k = 1, 2, ... , n — 1) about an axis of symmetry of the «th order is the element C n ~ k = C n n ~ k , i.e. a rotation through an angle (n—k)2irjn in the same direction or, what is the same thing, a rotation through an angle 2hn\n in the opposite direction. If, among the transformations in the group, there is a rotation through an angle it about a perpendicular axis (this rotation reverses the direction of the axis under consideration), then, by the general rule proved above, the rotations C n k and C n ~ k belong to the same class. A reflection a n in a plane perpendicular to the axis also reverses its direction ; however, it must be borne in mind that the reflection also changes the direction of rotation. Hence the existence of a h does not render C n k and C n ~ k conju- gate. A reflection <j v in a plane passing through the axis, on the other hand, does not change the direction of the axis, but changes the direction of rota- tion, and therefore C n ~ k = o v C n k cr v , so that C n k and C n ~ k belong to the same class if such a plane of symmetry exists. If rotations about an axis through the same angle in opposite directions are conjugate, we shall call it bilateral. The determination of the classes of a point group is often facilitated by the following rule. Let G be some group not containing the inversion /, and C i a group consisting of the two elements / and E. Then the direct product G x C i is a group containing twice as many elements as G ; half of them are the same as the elements of the group G, while the remainder are obtained by multiplying the latter by /. Since / commutes with any other transformation of a point group, it is clear that the group GxC t contains twice as many classes as G ; to each class A of the group G there correspond the two classes A and AI in the group G x C t . In particular, the inversion I always forms a class by itself. 340 The Theory of Symmetry §93 Let us now go on to enumerate all possible point groups. We shall con- struct these by starting from the simplest ones and adding new elements of symmetry. We shall denote point groups by bold italic Latin letters with appropriate suffixes. !• c n groups The simplest type of symmetry has a single axis of symmetry of the wth order. The group C n is the group of rotations about an axis of the nth order. This group is evidently cyclic. Each of its n elements forms a class by itself. The group C x contains only the identical transformation E, and corresponds to the absence of any symmetry. U- s zn groups The group S 2n is the group of rotary-reflections about a rotary-reflection axis of even order 2n. It contains 2n elements and is evidently cyclic. In particular, the group S 2 contains only two elements, E and /; it is also denoted by C^ We may note also that, if the order of a group is a number of the form 2n = 4^+2, inversion is among its elements; it is clear that (S 4p+2 ) 2p+1 = C 2 a h = I. Such a group can be written as a direct product S^ p+2 — C 2p+1 xC t ; it is also denoted by C 2p+1 4 . In » c nh groups These groups are obtained by adding to an axis of symmetry of the «th order a plane of symmetry perpendicular to it. The group C nh contains 2m elements: n rotations of the group C n and n rotary-reflection trans- formations C n k a h , k = 1, 2, ... , n (including the reflection C n n a h = a h ). All the elements of the group commute, i.e. it is Abelian; the number of classes is the same as the number of elements. If n is even (n = 2p), the group contains a centre of symmetry (since C 2p p a h = C 2 a h = /). The simplest group, C lh , contains only two elements, E and a h \ it is also denoted by C s . IV - c nv groups If we add to an axis of symmetry of the «th order a plane of symmetry passing through it, this automatically gives another n — 1 planes intersecting along the axis at angles of 7r/«, as follows at once from the geometrical theorem f (91.7) stated in §91. The group C nv thus obtained therefore con- tains 2« elements : n rotations about the axis of the «th order, and n reflections a v in vertical planes. Fig. 34 shows, as an example, the systems of axes and planes of symmetry for the groups C 3V and C iv . To determine the classes, we notice that, because of the presence of planes of symmetry passing through the axis, the latter is bilateral. The actual distribution of the elements among the classes depends on whether n is even or odd. f It is easy to see that, in a finite group, there cannot be two planes of symmetry intersecting at an angle which is not a rational fraction of 2w. If there were two such planes, it would follow that there were an infinite number of other planes of symmetry, intersecting along the same line and obtained by reflecting one plane in the other ad infinitum. In other words, if there are two such planes, there must be complete axial symmetry. §93 Point groups 341 If n is odd (n = 2/> + l), successive rotations C 2p+1 carry each of the planes successively into each of the other 2p planes, so that all the planes of symmetry are equivalent, and the reflections in them belong to a single class. Among rotations about the axis there are 2p operations apart from the identity, and these are conjugate in pairs, forming p classes each of two elements (C 2p +i fc and C 2p +r k , k = 1, 2, ... ,p); moreover, E forms an extra class. Thus there arep+2 classes altogether. X l/Sl ^ C 4¥ Fig. 34 If, on the other hand, n is even (n = 2p), only every alternate plane can be interchanged by successive rotations C 2p ; two adjacent planes cannot be carried into each other. Thus there are two sets of p equivalent planes, and accordingly two classes of p elements (reflections) each. Of the rotations about the axis, C 2p 2p = E and C 2p v = C 2 each form a class by themselves, while the remaining 2p— 2 rotations are conjugate in pairs and give another p— 1 classes, each of two elements. The group C 2p „ thus has />+3 classes altogether. V. D n groups If we add to an axis of symmetry of the nth order an axis of the second order perpendicular to it, this involves the appearance of a further n — 1 such axes, so that there are altogether n horizontal axes of the second order, intersecting at angles irjn. The resulting group D n contains 2n elements: n rotations about an axis of the nth order, and n rotations through an angle 77 about horizontal axes (we shall denote the latter by U 2 > reserving the notation C 2 for a rotation through an angle it about a vertical axis). Fig. 34 shows, as an example, the systems of axes for the groups D 3 and D 4 . In an exactly similar manner to case IV, we may verify that the axis of the nth order is bilateral, while the horizontal axes of the second order are all equivalent if n is odd, or form two non-equivalent sets if n is even. Con- sequently, the group D 2p has the following/) +3 classes: E, 2 classes each of p rotations U 2 , the rotation C 2 , and p — 1 classes each of two rotations about the vertical axis. The group D 2p+1 , on the other hand, has p+2 classes: E, 2p + 1 rotations C/ 2 , and p classes each of two rotations about the vertical axis. 342 The Theory of Symmetry §93 An important particular case is the group D 2 . Its system of axes is composed of three mutually perpendicular axes of the second order. This group is also denoted by V. VI - D nh groups If we add to the system of axes of a group D n a horizontal plane of sym- metry passing through the n axes of the second order, n vertical planes automatically appear, each of which passes through the vertical axis and one of the horizontal axes. The group D nh thus obtained contains 4« elements; besides the 2n elements of the group D n , it contains also n reflections a v and n rotary-reflection transformations C n k a h . Fig. 35 shows the system of axes and planes for the group D 3h . ^ h ^V 1/\J u 3h The reflection a h commutes with all the other elements of the group; hence we can write D nh as the direct product D nh = D n xC s , where C s is the group consisting of the two elements E and a h . For even n the inversion operation is among the elements of the group, and we can also write D 2j >,fc = Z> 23 ,xCV Hence it follows that the number of classes in the group D nh is twice the number in the group D n . Half of them are the same as those of the group D n (rotations about axes), while the remainder are obtained by multiplying these by a h . The reflections a v in vertical planes all belong to a single class (if n is odd) or form two classes (if n is even). The rotary-reflection trans- formations a h C n k and a h C n ~ k are conjugate in pairs. VI1 - D nd groups There is another way of adding planes of symmetry to the system of axes of the group D n . This is to draw vertical planes through the axis of the nth order, midway between each adjacent pair of horizontal axes of the second order. The adding of one such plane again involves the appearance of another (»— 1) planes. The system of axes and planes of symmetry thus obtained determines the group D nd . Fig. 35 shows the axes and planes for the groups D 2d andD 3d . §93 Point groups 343 The group D nd contains An elements. To the 2« elements of the group D n are added n reflections in the vertical planes (denoted by a d — the "diago- nal" planes) and n transformations of the form G = U 2 a d . In order to ascertain the nature of these latter, we notice that the rotation U 2 can, by (91.6), be written in the form U 2 = a h a v , where a v is a reflection in the verti- cal plane passing through the corresponding axis of the second order. Then G = a n a v a d (the transformations a v , a h alone are not, of course, among the elements of the group). Since the planes of the reflections a v and a d intersect along an axis of the wth order, forming an angle (2&+l)7r/2w, where &=l,...,(n— 1) (since here the angle between adjacent planes is tt/2w), it follows that, by (91.6), we have a v a d = C 2n 2k+1 . Thus we find that G = a h C 2n 2k+1 = S 2n 2k+1 , i.e. these elements are rotary-reflection trans- formations about the vertical axis, which is consequently not a simple axis of symmetry of the nth. order, but a rotary-reflection axis of the 2»th order. The diagonal planes reflect two adjacent horizontal axes of the second order into each other ; hence, in the groups under consideration, all axes of the second order are equivalent (for both even and odd n). Similarly, all diagonal planes are equivalent. The rotary-reflection transformations S2n 2k+1 and S2n~ 2k ~ 1 are conjugate in pairs.f Applying these considerations to the group D 2p d , we find that it contains the following 2p+3 classes: E, the rotation C 2 about the axis of the nth order, (p— 1) classes each of two conjugate rotations about the same axis, one class of the 2p rotations U 2 , one class of 2p reflections a d , and^ classes each of two rotary-reflection transformations. For odd n (= 2p+\), inversion is among the elements of the group; this is seen from the fact that, in this case, one of the horizontal axes is perpen- dicular to a vertical plane. Hence we can write D 2p+ld = D 2p+1 xC it so that the group D 2p+1 d contains 2p+4 classes, which are obtained at once from the p+2 classes of the group D 2p+1 . VIII. The group T (the tetrahedron group) The system of axes of this group is the system of axes of symmetry of a tetrahedron. It can be obtained by adding to the system of axes of the group V four oblique axes of the third order, rotations about which carry the three axes of the second order into one another. This system of axes is conveniently represented by showing the three axes of the second order as passing through the centres of opposite faces of a cube, and those of the third order as the spatial diagonals of the cube. Fig. 36 shows the position of these axes in a cube and in a tetrahedron (one axis of each type is shown). The three axes of the second order are mutually equivalent. The axes of the third order are also equivalent, since they are carried into one another by f For we have "A* 1 ** 1 ** = WV** 1 ** = WV*«<r, = a/An" 2 *- 1 = s 2n - 2 *-i 344 The Theory of Symmetry §93 Fig. 36 the rotations C 2 , but they are not bilateral axes. Hence it follows that the twelve elements in the group T are divided into four classes : E, the three rotations C 2 , the four rotations C 3 and the four rotations C 3 2 . IX. The group T d This group contains all the symmetry transformations of the tetrahedron. Its system of axes can be obtained by adding to the axes of the group T planes of symmetry, each of which passes through one axis of the second order and two of the third order. The axes of the second order thereby become rotary- reflection axes of the fourth order (as in the case of the group D 2d ). This system is conveniently represented by showing the three rotary-reflection axes as passing through the centres of opposite faces of a cube, the four axes of the third order as its spatial diagonals, and the six planes of symmetry §93 Point groups 345 as passing through each pair of opposite edges (Fig. 37 shows one of each kind of axis and one plane). Since the planes of symmetry are vertical with respect to the axes of the third order, the latter are bilateral axes. All the axes and planes of a given kind are equivalent. Hence the 24 elements of this group are divided into the following five classes: E, eight rotations C 3 and C 3 2 , six reflections in planes, six rotary-reflection transformations 5 4 and 5 4 3 , and three rotations C — S 2 X. The group T h This group is obtained from T by adding a centre of symmetry: T h = TxCp As a result, three mutually perpendicular planes of symmetry appear, passing through each pair of axes of the second order, and the axes of the third order become rotary-reflection axes of the sixth order (Fig. 38 shows one of each kind of axis and one plane). / ' A 'A (~\- K n__ 7 7 4-^ V-j,' Fig. 38 The group contains 24 elements divided among eight classes, which are obtained at once from those of the group T. XI. The group O (the octahedron group) The system of axes of this group is the system of axes of symmetry of a cube: three axes of the fourth order pass through the centres of opposite 346 The Theory of Symmetry §93 faces, four axes of the third order through opposite corners, and six axes of the second order through the midpoints of opposite edges (Fig. 39). It is easy to see that all the axes of a given order are equivalent, and each of them is bilateral. Hence the 24 elements are divided among the following five classes: E, eight rotations C 3 and C 3 2 , six rotations C 4 and C 4 3 , three rotations C 4 2 and six rotations C 2 . XII. The group O h This is the group of all symmetry transformations of the cube.f It is obtained by adding to the group O a centre of symmetry: O h = OxC^ The axes of the third order in the group O are thereby converted into rotary- reflection axes of the sixth order (the spatial diagonals of the cube); in addition, another six planes of symmetry appear, passing through each pair of opposite edges, and three planes parallel to the faces of the cube (Fig. 40). The group contains 48 elements divided among ten classes, which Fig. 40 can be at once obtained from those of the group O ; five classes are the same as those of the group O, while the remainder are : /, eight rotary-reflection transformations S 6 and 5 6 5 , six rotary-reflection transformations C^a^ C 4 3 a h about axes of the fourth order, three reflections a h in planes horizontal with respect to the axes of the fourth order, and six reflections a v in planes vertical with respect to these axes. XIII, XIV. The groups Y, Y h {the icosahedron groups) These groups are of no physical interest, since they do not occur in Nature as symmetry groups of molecules. Hence we shall here only mention that Y is a group of 60 rotations about the axes of symmetry of the icosa- hedron (a regular solid with twenty triangular faces) or of the pentagonal dodecahedron (a regular solid with twelve pentagonal faces); there are six axes of the fifth order, ten of the third and fifteen of the second. The group Y h is obtained by adding a centre of symmetry: Y h = YxC it and is the f The groups T, T d , T h , O, 0% are often called cubic. §94 Representations of groups 347 complete group of symmetry transformations of the above-mentioned poly- hedra. This exhausts all possible types of point group containing a finite number of elements. In addition, we must consider what are called continuous point groups, which contain an infinite number of elements. This we shall do in §98. §94. Representations of groups Let us consider any symmetry group, and let fa be some one-valued func- tion of the co-ordinates."}* Under the transformation of the co-ordinate system which corresponds to an element G of the group, this function is changed into some other function. On performing in turn all the g transfor- mations in the group {g being the order of the group), we in general obtain g different functions from fa. For certain fa, however, some of these functions may be linearly dependent. As a result we obtain some number f(<g) of linearly independent functions fa, fa, ... , ifj f , which are transformed into linear combinations of one another under the transformations belonging to the group in question. In other words, as a result of the transformation G, each of the functions fa (i = 1, 2, 3, ... ,/) is changed into a linear combina- tion of the form where the G ik are constants depending on the transformation G. The array of these constants is called the matrix of the transformation. J In this connection it is convenient to regard the elements G of the group as operators acting on the functions tft^ so that we can write * = S«; (94.1) the functions fa can always be chosen so as to be orthonormal. Then the concept of the matrix of the transformation is the same as that of the matrix of the operator, in the form defined in §11: G ik =jfa*Gfadq. (94.2) To the product of two elements G and H of the group there corresponds the matrix obtained from the matrices of G and H by the ordinary rule of matrix multiplication (11.12): {GH) ik = ZG a H lk . (94.3) The set of matrices of all the elements in a group is called a representation t In the configuration space of the physical system concerned. t Since the functions i[n are assumed one-valued, a definite matrix corresponds to each element of the group. 348 The Theory of Symmetry §94 of the group. The functions tf/ 1} ... , tf/ f with respect to which these matrices are defined are called the basis of the representation. The number/ of these functions gives what is called the dimension of the representation. Let us consider the integral j \ip\ 2 dq, where \\s is some function of the co-ordinates. Since the integral is taken over all space, it is evident that its value is unchanged by any rotation or reflection of the co-ordinate system. Hence, for any symmetry transformation G, we can write f (0*<P*)(OiJj) dq = f 0*0 dq. Introducing the transposed operator 0, we have f (0*ift*)(Otf>) dq = ! 0(5(5*0* dq = ! 0*0 dq, whence, from the fact that is arbitrary, it follows that 00* = 1, or 0* = 0~\ i.e., the operators are unitary. Thus a representation of a symmetry group in terms of orthonormal base functions is unitary, i.e. the group is represented by unitary matrices. Suppose that we perform on the system of functions 0i, . . . , 0/ the linear unitary transformation 0' f =£0*. (94.4) This gives a new system of functions 0'i, . . . , 0'/, which are also orthonormal (see §12).| If we now take, as the basis of the representation, the functions ifj'i, we obtain a new representation of the same dimension. Such representa- tions, obtained from one another by a linear transformation of their base functions, are said to be equivalent ; it is evident that they are not essentially different. The matrices of equivalent representations can be simply expressed in terms of one another. According to (12.7), the matrix of the operator in the new representation is the matrix of the operator 0' =S-iO§ (94.5) in the old representation. The sum of the diagonal elements (i.e. the trace) of the matrix representing an element G of a group is called its character; we shall denote it by x(G). It is a very important result that the characters of the matrices of equivalent f It may also be noted that the unitary transformation leaves invariant the sum of the squared moduli of the functions: using (12.6), we have E|0' i\ 2 = J ZSk t faS li fp = S^i0i«8jM = S|$t| 2 . §94 Representations of groups 349 representations are the same (see the end of §12). This circumstance gives particular importance to the description of group representations by stating their characters: it enables us to distinguish at once the fundamentally different representations from those which are equivalent. Henceforward we shall regard as different representations only those which are not equivalent. If we take 5 in (94.5) to be that element of the group which relates the conjugate elements G and G', we have the result that, in any given represen- tation of a group, the characters of the matrices representing elements of the same class are the same. The identical transformation corresponds to the unit element E of the group. Hence the matrix representing the latter is diagonal in every represen- tation, and the diagonal elements are unity. The character x( E ) is con- sequently just the dimension of the representation : X(E) =/• (94.6) Let us consider some representation of dimension/. It may happen that, as a result of a suitable linear transformation (94.4), the base functions divide into sets of f v f 2 , ... functions (/i4-/ a +- =/)> in such a way that, when any element of the group acts on them, the functions in each set are transformed only into combinations of themselves, and do not involve functions from other sets. In such a case the representation in question is said to be reducible. If, on the other hand, the number of base functions that are transformed only into combinations of themselves cannot be reduced by any linear trans- formation of them, the representation which they give is said to be irreducible. Any reducible representation can, as we say, be decomposed into irreducible ones. This means that, by the appropriate linear transformation, the base functions divide into several sets, of which each is transformed by some irreducible representation when the elements of the group act on it. Here it may be found that several different sets transform by the same irreducible representation; in such a case this irreducible representation is said to be contained so many times in the reducible one. Irreducible representations are an important characteristic of a group, and play a fundamental part in all quantum-mechanical applications of group theory. We shall give the chief properties of irreducible representations.f It may be shown that the number of different irreducible representations of a group is equal to the number r of classes in the group. We shall distin- guish the characters of the various irreducible representations by indices ; the characters of the matrices of the element G in the representations are x d)(G), X < 2 >(G), ... , X <->(G). t The proof of these properties may be found in books on group theory, for example E. Wigner, Group Theory and its Applications to the Quantum Mechanics of Atomic Spectra, Academic Press, New York 1959 (Gruppentheorie und ihre Anwendung auf die Quantenmechanik der Atomspektren, Vieweg, Brunswick 1931); H. Weyl, The Theory of Groups and Quantum Mechanics, Methuen, London 1931. 350 The Theory of Symmetry §94 The matrix elements of irreducible representations satisfy a number of ortho- gonality relations. First of all, for two different irreducible representations the relations SGW tt GW> Im *=0 (94.7) hold, where a and /J (a # j8) refer to the two irreducible representations, and the summation is taken over all the elements of the group. For any irreducible representation the relations |GW=^A,» (94.8) Ja. hold, i.e. only the sums of the squared moduli of the matrix elements are not zero: §|G<«y 2 =£// a . The relations (94.7), (94.8) can be combined in the form S &«\ k GV» lm * = fs^Am- (94.9) JOL In particular, we can obtain from this an important orthogonality relation for the characters of the representations. Summing both sides of equation (94.9) over equal values of the suffixes i, k and /, m, we have S x («)(G) x ^(G)*=^S a/? . (94.10) For a = j8 we have glx (a, (G)l 2 =£, i.e. the sum of the squared moduli of the characters of an irreducible represen- tation is equal to the order of the group. We may notice that this relation can be used as a criterion of the irreducibility of a representation; for a reducible representation, this sum is always greater than g (it is ng, where n is the number of irreducible representations contained in the reducible one). It also follows from (94.10) that the equality of the characters of two irreducible representations is not only a necessary but also a sufficient con- dition for them to be equivalent. Since the characters of elements of the same class are equal, the sum (94.10) actually contains only r independent terms, and can be written in the form ^gcx^{C) X ^{C)*=go^ (94.11) §94 Representations of groups 351 where the summation is over the r classes of the group (arbitrarily denoted by C) and^c is the number of elements in class C.f The relation (94.10) enables any reducible representation to be very easily decomposed into irreducible ones if the characters of both are known. Let x(G) be the characters of some reducible representation of dimension /, and let the numbers a (1) , a {2) , ... , a {r) indicate how many times the corres- ponding irreducible representations are contained in it, so that £ d% =/, (94.12) p-i where fp are the dimensions of the irreducible representations. Then the characters x(G) can be written X (G) = 2 a^x^(G). (94.13) p -l Multiplying this equation by x (a) (^)* an & summing over all G, we have by (94.10) i a*">=-Sx(G) x («>(G)*. (94.14) § ° Let us consider a representation of dimension f — g, given by the g functions 0$, tft being some general function of the co-ordinates (so that all the g functions Oi/j obtained from it are linearly independent); such a representation is said to be regular. It is clear that none of the matrices of this representation will contain any diagonal elements, with the exception of the matrix corresponding to the unit element ; hence x(G) = for G =£ E, while x{E) = S- Decomposing this representation into irreducible ones, we have for the numbers a^\ by (94.14), the values a™ = (llg)gf (0L) =/ (a) , i.e. each irreducible representation is contained in the reducible one under consideration as many times as its dimension. Substituting this in (94.12), we find the relation /i 2 +/ 2 2 +...+/ r 2 =iT, (94-15) the sum of the squared dimensions of the irreducible representations of a group is equal to its order .J Hence it follows, in particular, that for Abelian groups (where r = g) all the irreducible representations are of dimension one (/i=/ a = ..•= /,= !)• t Since the number of irreducible representations is equal to the number of classes, the quantities f ac = ■\/( < gclg)x^XC) form a square matrix of r 2 quantities. The orthogonality relations for the first suffix (2/ac//3c* = Sa/3) then automatically give those for the second suffix: S/ac/ac'* = hcc'. Hence, besides (94.11), we have Sx (a) (C)x (a) (C")* = — Sec-. (94.11a) a gc % It may be mentioned that, for point groups, equation (94.15) for given r and g can in practice be satisfied in only one way by a set of integers / 1( ... ,/ r . 352 The Theory of Symmetry §94 We may also remark, without proof, that the dimensions of the irreducible representations of a group divide its order. Among the irreducible representations of any group there is always a trivial one, given by a single base function invariant under all the trans- formations in the group. This one-dimensional representation is called the unit representation ; in it, all the characters are unity. Let us consider two different systems of functions $t (a) , ... , if) f (a) and ti^K ... , "A/g^, which form two irreducible representations of a group. By forming the products «A < (a) «/'jfc (/?) we obtain a system off a fp new functions, which can serve as the basis for a new representation of dimension f u fp. This representation is called the direct product of the other two; it is irreduc- ible only if f a orfp is unity. It is easy to see that the characters of the direct product are equal to the products of the characters of the two component representations. For, if I m then hence we have for the characters, which we denote by (x^Xx^X^D* (x (a) Xx^)(<?) = xgjcog^ = 2 G w s G tt (0, i.e. ( x «*) Xx (/?))(G) = x m (G)x^(G). (94.16) The two irreducible representations so multiplied may, in particular, be the same; in this case we have two different sets of functions if/ ly ... , iftf and <f) x , ... , <f) f giving the same representation, while the direct product of the representation with itself is given by the f 2 functions ifti<f> k , and has the characters (xXxXG) = b((G)] 2 - This reducible representation can be at once decomposed into two represen- tations of smaller dimension (although these are, in general, themselves reducible). One of them is given by the i/(/+l) functions ^&+^a;<^> the other by the \f{f— 1) functions 0^&— «Afc<^, i ¥=■ k; it is evident that the func- tions in each of these sets are transformed only into combinations of them- selves. The former is called the symmetric product of the representation with itself, and its characters are denoted by the symbol [x 2 J(G);the latter is called the antisymmetric product, and its characters are denoted by {x 2 } (G). To determine the characters of the symmetric product, we write G(rpi<f> k +*p k $i) =£ G M G r ?nfc (j/f,^ m +^ tn ^ J ) = I ^(GuGrnk+GmiGt^ifiicfin+tfimfa). §94 Representations of groups 353 Hence we have for the character [X 2 ](G) = \ §(<?«<?»+<?*<?«). But S(? w =x(G), and £ G ik G ki = x(G 2 ) ; thus we finally obtain the formula \X*](G) = ttbc(G)?+x(G*)}> (94.17) which enables us to determine the characters of the symmetric product of a representation with itself from the characters of the representation. In an exactly similar manner, we find for the characters of the antisymmetric product the formula (X 2 }(G) = m(G)]*-x(G*)}. (94.18) If the functions tft i and <^ are the same, we can evidently construct from them only the symmetric product, formed by the squares ^ 2 and the pro- ducts iff^k, i ¥" k. In applications, symmetric products of higher orders are also encountered; their characters may be obtained in a similar manner. An important property of direct products is the following. The resolution of the direct product of two different irreducible representations into irre- ducible parts never involves the unit representation, but the direct product of an irreducible representation with itself (that is, of course, its symmetric part) always contains (and only once) the unit representation. In order to know whether the unit representation is present in the representation (94.16), we simply sum its characters with respect to G and divide the result by the order g of the group, in accordance with (94.14). From (94.10) it is seen that this gives zero if a ^ j3 and unity if a = j8. For applications it is useful to know a formula which enables us to represent an arbitrary function iff as a sum of functions transformed by the irreducible representations of the group, i.e. in the form = SE&< a >, (94.19) where the functions «/f i (a) (i = 1, 2, ... ,/ a ) are transformed by the ath irreducible representation. The problem consists in determining the func- tions < < a > from the function tft, and is solved by the formula rfjfoo = {5L 2 Gi^Gj*. (94.20) 8 G To prove this, it suffices to show that the expression on the right-hand side of the equation reduces to ^r 4 (a) identically if we put ijj = ^ (a) , and to zero if we put tft = ip k W with k # i or j8 # a; both these results follow at once from the orthogonality relations (94.7), (94.8), putting G$ k W = S G lk ^^\ 354 The Theory of Symmetry §95 If we substitute (94.20) in (94.19) and effect the summation over t, we obtain a simpler expansion of the arbitrary function xjj into functions ^t (a) belonging to the various irreducible representations, but not to any definite rows of these representations : = 2 «A (a) , </-«*> = — S x < a >(G)*&A. (94.21) Finally, we shall make a few remarks regarding the irreducible represen- tations of a group which is the direct product of two other groups (not to be confused with the direct product of two representations of the same group). If the functions «/r f (a) give an irreducible representation of the group A, and the functions <f> k W give one of the group B, the products ^ A (/?) ^ f (ot) are the basis of an / a /^-dimensional representation of the group A xB, and this representation is irreducible. The characters of this representation are obtained by multiplying the corresponding characters of the original represen- tations (cf. the derivation of formula (94.16)); to an element C= AB of the group A xB there corresponds the character X(Q= X M(A) X W(B). (94.22) Multiplying together in this way all the irreducible representations of the groups A and B, we obtain all the irreducible representations of the group AxB. §95. Irreducible representations of point groups Let us pass now to the actual determination of the irreducible represen- tations of those point groups which are of physical interest. The great majority of molecules have axes of symmetry only of the second, third, fourth or sixth order. Hence it is unnecessary to consider the icosahedron groups Y, Y h ; we shall examine the groups C n , C nhi C nv , D n , D nh only for the values n = 1, 2, 3, 4, 6, and the groups S 2ny D nd only for n = 1, 2, 3. The characters of the representations of these groups are shown in Table 7. Isomorphous groups have the same representations and are given together. The numbers in front of the symbols for the elements of a group in the upper rows show the numbers of elements in the corresponding classes (see §93). The left-hand columns show the conventional names usually given to the representations. The one- dimensional representations are denoted by the letters A, B, the two-dimensional ones by E, and the three-dimensional ones by F; the notation E for a two-dimensional irreducible representation should not be confused with the unit element of a group.f The base functions of A representations are symmetric, and those of B representations antisym- metric, with respect to rotations about a principal axis of the nth. order. t The reason why two complex conjugate one-dimensional representations are shown as one two-dimensional one is explained in §96. §95 Irreducible representations of point groups 355 The functions of different symmetry with respect to a reflection an are distinguished by the number of primes (one or two), while the suffixes g and u show the symmetry with respect to inversion. Beside the symbols for the representations are placed the letters x, y, z; these show the repre- sentations by which the co-ordinates themselves are transformed (with a view to later quantum-mechanical applications). The z-axis is always taken along Table 7 Characters of irreducible representations of point groups c x E c t c 2 C s E E E c 2 a c 3 E C 3 C 8 A 1 A\z 1 1 1 E;x±iy j 1 1 e e 2 A g A;z A';x,y 1 1 e 2 e Au\ x,y,z B;x,y A"\z 1 -1 (-■2h E c 2 Oh / c 2e E c 2 a v O'v V^D 2 E c 2 * c 2 y c 2 * Ag A l ;z A 1 1 l 1 Bg B 2 ;y B 3 ;x 1 -1 -l 1 Au\ z A 2 B t ;z 1 1 -l -1 Bu\ X, y B x ;x B 2 ;y 1 -1 l -1 ^3« E 2C S 3<7„ D z E 2C S 3C/ 2 A t ;z A 1 1 1 1 A 2 A 2 \ z 1 1 -1 E;x, y E;x,y 2 -1 c. E d C 2 C 4 S s t E S 4 C 2 S 4 * A\z A 1111 B B;z 1-11 -1 E;x±iy E;x±iy \ 1 i -1 -i 1 -i -1 i C t E C 6 c 3 c 2 C 3 2 cv A;z 1 1 1 1 1 B -1 1 -1 1 -1 f O) 2 — m 1 CO 2 — CO E 1 I — a> u,* 1 — CO CO 2 CO CO 2 -1 — CO -co 2 E 2 ; x ±iy 1 -to 2 — CO -1 to* CO 356 The Theory of Symmetry Table 7 — continued §95 c 4c E c 2 2C 4 2<7„ 2a' \ D t E c % 2C 4 2U 2 2U\ D 2d E c 2 2S t 2U 2 2a a A x \z A t A 1 1 1 1 1 1 A 2 A 2 \ z A 2 1 1 1 -1 -1 B 1 B 1 B 1 1 1 -1 1 -1 B 2 B 2 B 2 ;z 1 1 -1 -1 1 E;x,y E;x,y E;x,y 2 -2 D* E c 2 2C 8 2C e 3U 2 3U' 2 ^60 E c 2 2C 8 2C 6 3ff„ 3ct'„ D ah E Oh 2C 8 2S 3 3U 2 3cr'„ A x Aaz AS 1 1 1 1 1 1 A 2 \z A % A,' 1 1 1 1 -1 -1 B 1 B 2 A," 1 -1 1 -1 1 -1 B 2 B, A 2 ; z 1 -1 1 -1 -1 1 E 2 E 2 E';x y 2 2 -1 -1 E x ;x,y Ei;x,y E" 2 -2 -1 1 o E 8C 3 3C 2 6C 2 6Ct T E 3C a 4C 8 4C 3 2 T d E 8C 8 3C 2 6a,i 6S^ A 1 1 1 1 A, A 1 1 1 1 1 1 * { 1 1 e e* A 2 A 2 1 1 1 -1 -1 I 1 1 e* e E E 2 -1 2 F; x, y, z 3 -1 F 2 Ftix F 2 ;x, y,z Fj. y,z 3 3 -1 1 -1 -1 -1 1 §95 Irreducible representations of point groups 357 the principal axis of symmetry. The letters e and a> denote e = «**/», to = e 2ffi/6 = -co 4 . The simplest problem is to determine the irreducible representations for the cyclic groups (C n , S n ). A cyclic group, like any Abelian group, has only one-dimensional representations. Let G be a generating element of the group (i.e. one which, on being raised to successive powers, gives all the elements of the group). Since G e = E (where g is the order of the group), it is clear that, when the operator acts on a base function ip, the latter can be multi- plied only by l 1 /*, i.e.f Cty^ftvaf (A = 1,2, ...,£). The group C 2 n (and the isomorphous groups C 2v and D 2 ) is Abelian, so that all its irreducible representations are one- dimensional, and the characters can only be ± 1 (since the square of every element is E). Next we consider the group C 3V . As compared with the group C 3 , the reflections a v in vertical planes (all belonging to one class) are here added. A function invariant with respect to rotation about the axis (a base function of the representation A of the group C 3 ) may be either symmetric or anti- symmetric with respect to the reflections a v . Functions multiplied by € and e 2 under the rotation C 3 , on the other hand (base functions of the com- plex conjugate representations £), change into each other on reflection. J It follows from these considerations that the group C sv (and D 3 , which is isomorphous with it) has two one-dimensional irreducible representations and one two-dimensional, with the characters shown in the table. The fact that we have indeed found all the irreducible representations may be seen from the result l 2 + l 2 +2 2 = 6, which is the order of the group. Similar considerations give the characters of the representations of other groups of the same type (C 4 „, C 6V ). The group T is obtained from the group V by adding rotations about four oblique axes of the third order. A function invariant with respect to transformations of the group V (a basis of the representation A) can be multiplied, under the rotation C 3 , by 1, e or e 2 . The base functions of the three one- dimensional representations B ly B 2 , Bz of the group V change into one another under rotations about the axis of the third order (this is seen, for example, if we take as these functions the co-ordinates x, y, z themselves). Thus we obtain three one-dimensional irreducible representations and one three-dimensional (l 2 + l 2 + l 2 + 3 2 = 12). Finally, let us consider the isomorphous groups O and T d . The group T d is obtained from the group T by adding reflections a d in planes each of Which passes through two axes of the third order. A base function of the f For the point group C n we can, for example, take as the functions ift the functions e tk *, k = 1, 2, ... ,n, where <f> is the angle of rotation about the axis, measured from some fixed direction. ' % These functions may, for example, be taken as fa == c'*, fa = e~i<l>. On reflection in a vertical plane, <f> changes sign. 358 The Theory of Symmetry §96 unit representation A of the group T may be symmetric or antisymmetric with respect to these reflections (which all belong to one class), and this gives two one- dimensional representations of the group T d . Functions multiplied by e or e 2 under a rotation about an axis of the third order (the basis of the complex conjugate representations E of the group T) change into each other on reflection in a plane passing through this axis, so that one two-dimensional representation is obtained. Finally, of three base functions of the representation F of the group T, one is transformed into itself on reflection (and can either remain unaltered or change sign), while the other two change into each other. Thus we have altogether two one-dimensional representations, one two-dimensional and two three-dimensional. The representations of the remaining point groups in which we are inter- ested can be obtained immediately from those already given, if we notice that the remaining groups are direct products of those already considered with the group C i (or C s ) : ^3h = C'sXCg Ceh = CeXC,- D 2h = D 2 xC € D, h = D 4 xQ "6 = ^3XC t - D 3d = D^xCt D 6h = D 6 xC f T h = TxC t o h =OxCi Each of these direct products has twice as many irreducible representations as the original group, half of them being symmetric (denoted by the suffix g) and the other half antisymmetric (suffix u) with respect to inversion. The characters of these representations are obtained from those of the representa- tions of the original group by multiplying by + 1 (in accordance with the rule (94.22)). Thus, for instance, we have for the group Dza the repre- sentations : D 3d E 2C 3 3*7 2 I 25 6 3(7,, A Xo 1 1 1 1 A 2 g 1 -1 1 -1 K 2 — 1 2 — 1 ■A-iu 1 1 -1 — 1 -1 A 2u 1 -1 -1 — 1 1 E u 2 — 1 -2 §96. Irreducible representations and the classification of terms The quantum-mechanical applications of group theory are based on the fact that the Schrodinger's equation for a physical system (an atom or §96 Irreducible representations and the classification of terms 359 molecule) is invariant with respect to symmetry transformations of the system. It follows at once from this that, on applying the elements of a group to a function satisfying Schrodinger's equation for some value of the energy (an eigenvalue), we must again obtain solutions of the same equation for the same value of the energy. In other words, under a symmetry transformation the wave functions of the stationary states of the system belonging to a given energy level transform into linear combinations of one another, i.e. they give some representation of the group. An important fact is that this representa- tion is irreducible. For functions which are invariably transformed into linear combinations of themselves under symmetry transformations must belong to the same energy level ; the equality of the eigenvalues of the energy cor- responding to several groups of functions (into which the basis of a reducible representation can be divided), which are not transformed into combinations of one another, would be an utterly improbable coincidence (provided that there is no special reason for such equality ; see below). Thus, to each energy level of the system, there corresponds some irreduc- ible representation of its symmetry group. The dimension of this represen- tation determines the degree of degeneracy of the level concerned, i.e. the number of different states with the energy in question. The fixing of the irreducible representation determines all the symmetry properties of the given state, i.e. its behaviour with respect to the various symmetry trans- formations. Irreducible representations of dimension greater than one are found only in groups containing non-commuting elements; Abelian groups have only one- dimensional irreducible representations. It is apposite to recall here that the relation between degeneracy and the presence of operators which do not commute with one another (but do commute with the Hamiltonian) has already been found above from considerations unrelated to group theory (§10). The following important reservation should be made regarding all these statements. As has already been pointed out (§18), the symmetry (valid in the absence of a magnetic field) with respect to a change in the sign of the time has, in quantum mechanics, the result that complex conjugate wave functions must belong to the same eigenvalue of the energy. Hence it follows that, if some set of functions and the set of complex conjugate functions give different irreducible representations of a group, these two complex conjugate representations must be regarded, from the physical point of view, as forming together a single representation of twice the dimension. In the preceding section we had examples of such representations. Thus the group C 3 has only one-dimensional representations; however, two of these are complex conjugates, and correspond physically to doubly degenerate energy levels. (In the presence of a magnetic field there is no symmetry with respect to a change in the sign of the time, and hence complex conjugate representations correspond to different energy levels.) Let us suppose that a physical system is subjected to the action of some 360 The Theory of Symmetry §96 perturbation (i.e. the system is placed in an external field). The question arises to what extent the perturbation can result in a splitting of the degener- ate levels. The external field has itself a certain symmetry.-}- If this symmetry is the same as or higher J than that of the unperturbed system, the symmetry of the perturbed Hamiltonian & = fi + f is the same as the symmetry of the unperturbed operator fi . It is clear that, in this case, no splitting of the degenerate levels occurs. If, however, the symmetry of the pertur- bation is lower than that of the unperturbed system, then the symmetry of the Hamiltonian i? is the same as that of the perturbation V. The wave functions which gave an irreducible representation of the symmetry group of the operator i? will also give a representation of the symmetry group of the perturbed operator #, but this representation may be reducible, and this means that the degenerate level is split. We shall show by means of an example how the mathematical techniques of group theory enable us to solve the problem of the splitting of any given level. Let the unperturbed system have symmetry Ta, and let us consider a triply degenerate level corresponding to the irreducible representation F 2 of this group. The characters of this representation are E 8C 3 3C 2 6a d 6S4 3 0-11-1 Let us assume that the system is subjected to the action of a perturbation with symmetry Cs v (with the third-order axis coinciding with one of those of the group Ta). The three wave functions of the degenerate level give a representation of the group Cz v (which is a sub-group of the group T a ), and the characters of this representation are equal to those of the same elements in the original representation of the group Ta, i.e. E 2C3 3g v This representation, however, is reducible. Knowing the characters of the irreducible representations of the group Czv, it is easy to decompose it into irreducible parts, using the general rule (94.14). Thus we find that it consists of the representations A\ and E of the group C% v . The triply degenerate level F2 is therefore split into one non- degenerate level A\ and one doubly degenerate level E. If the same system is subjected to the action of a per- turbation of symmetry Czv, which is also a sub-group of the group Ta, then t For example, in the case of the energy levels of the d and/ shells of ions in a crystal lattice which interact slightly with the surrounding atoms, the perturbation (the external field) is the field acting on an ion due to the other atoms. % If a symmetry group H is a sub-group of the group G, we say that H corresponds to a lower symmetry and Gtoa higher symmetry. It is evident that the symmetry of the sum of two expressions, one of which has the symmetry of G and the other that of H, is the lower symmetry, that of H. §97 Selection rules for matrix elements 361 the wave functions of the same level F 2 give a representation with characters 3 ~i i r Decomposing this into irreducible parts, we find that it contains the repre- sentations Ai, Bi, B 2 . Thus in this case the level is completely split into three non- degenerate levels. §97. Selection rules for matrix elements Group theory not only enables us to carry out a classification of the terms of any symmetrical physical system, but also gives us a simple method of finding the selection rules for the matrix elements of the various quantities which characterise the system. This method is based on the following general theorem. Let «/»/«> be one of the base functions of an irreducible (non-unit) representation of a symmetry group. Then the integral of this function over all spacef vanishes identically : f ,/,.(«) d? =0. (97.1) The proof is based on the evident fact that the integral over all space is invariant with respect to any transformation of the co-ordinate system, including any symmetry transformation. Hence J #* dq = J &A/ a > dq = J* S Gtpty,?* dq. We sum this equation over all the elements of the group. The integral on the left is simply multiplied by g, the order of the group, and we have g^Wdq-ZJ^ZG^dq. However, for any non-unit irreducible representation we have identically S G k M = 0; G k% this is a particular case of the orthogonality relations (94.7), when one of the irreducible representations is the unit representation. This proves the theorem. If j/r is a function belonging to some reducible representation of a group, the integral J «/» dq will be zero except when this representation contains the unit representation. This theorem is a direct consequence of the previous one. t That is, the configuration space of the physical system concerned. 362 The Theory of Symmetry §97 Let / be the operator of some scalar physical quantity. By definition, it is invariant with respect to all symmetry transformations. Its matrix elements are the integrals j *f,M*fy k W dq, (97.2) where the indices a, j8 distinguish different terms of the system, and the suffixes j, k denumerate the wave functions of states belonging to the same degenerate term. We denote the irreducible representations of the symmetry group of the system concerned that are given by the functions >> and if/ k W by the symbols D™ and W>. The products faM fc W give the representa- tion D<«>xZ)W; since the operator /itself is invariant with respect to all transformations, the whole expression in the integrand belongs to this representation. The direct product of two different irreducible represen- tations, however, does not contain the unit representation (see §94), whilst the direct product of an irreducible representation with itself always contains the unit representation, and only once; the integrals (97.2) are a constant (independent of i and k) times 8 ik . Thus we reach the conclusion that, for a scalar quantity, the matrix elements are non-zero only for transitions between states of the same type (i.e. belonging to the same irreducible representation). This is the most general form of a theorem of which we have already met several particular cases. Let us next consider some vector physical quantity A. The three compo- nents A x , A y , A z transform into linear combinations of themselves under symmetry transformations, and therefore give some representation of the symmetry group, which we shall denote by| D A . The products «A i (a) A^ & ( ^ ) give the representation D < - a) xD A xD^ ) ; the matrix elements are non-zero if this representation contains the unit representation. In practice, it is more convenient to decompose into irreducible parts the direct product Z) (os) x D A ; this gives us immediately all the types Z) W of states for transitions into which (from a state of type D (a) ) the matrix elements are not zero. The diagonal matrix elements (unlike those for transitions between diff- erent states of the same type) require special consideration. In this case we have only one system of functions ^//>>, not two different ones, and their products in pairs give the symmetric product [Z) (a > 2 ] of the representation Z> (a) with itself, not the direct product Z>( a >xD< a >. Hence the presence of diagonal matrix elements of a vector quantity means that the unit represen- tation is present in the decomposition of the product [Z) (a)2 ] XD A , or, what is the same thing, that D A is present in [Z) (a)2 ]|. Similarly we can find the selection rules for the matrix elements of a tensor. Examples of the application of these rules are given in the following Problems. f In general Da is different for polar and axial vectors. J We did not make this remark in considering the scalar quantity/, seeing that the symmetric pro- duct [D*") 2 ], like the direct product Z)( a >xZ)< a ), always contains the unit representation. Hence the diagonal matrix elements of a scalar quantity are, in general, different from zero. §97 Selection rules for matrix elements 363 PROBLEMS Problem 1. Find the selection rules for the matrix elements of a polar vector when symmetry O is present. Solution. The components of a vector are transformed by the irreducible representation F x . The decompositions of the direct products of F x with the other representations of the group O are F x xA x = F ly F ± xA 2 = F 2 , F ± xE = F x +F 2 , F x xF x = A x +E+F x +F ty F x xF 2 = A^E+F^F^ (1) Hence the non-zero non-diagonal matrix elements are those for the transitions F 1 <^. A lf E, F ly F 2 , F 2 4-> A 2 , E, F 2 . The symmetric products of the irreducible representations of the group O are W] = W] = A, [E 2 ] = A x +E, [*?] = [Ff] = A x +E+F v (2) F x is contained in none of these; hence the diagonal matrix elements vanish. Problem 2. The same as Problem 1, but for symmetry Z> 3 <f. Solution. The ^-component of the vector is transformed by the representation Am, the * and y components by E u - We have E u xA lg = E u xA 2g = E u , E U XA 1U = E u xA 2u = E g , E u xE u = A lg +A 2g +E g , E u xE g = A lu +A 2u +E u . Hence the non-diagonal matrix elements oiA x ,A v are non-zero for the transitions E u <— > A ig , A 29 , Eg',E g <-+ A 1U , A 2U . In the same way we find the selection rules for the matrix elements of At : A ig <-> A w , A 2B <-» A 1U , Eg <-> E u . The symmetric products of the irreducible representations are [A 1B 2 ] = [Au 2 ] = W] = [AJ] = A u , [E g 2 ] = [E u *] = E g +A lg . These do not contain either A 2U or E u \ hence the diagonal matrix elements vanish for both At and A x , A y . Problem 3. Find the selection rules for the matrix elements of a symmetrical tensor A ik of rank two (with Axx+A yy +Azz = 0) when symmetry O is present. Solution. The components A xy , A X z, A yz are transformed by F % . Decomposing the direct products of F 2 with all the representations of the group O, we find the selection rules F l < r ^A i , E, F lt F 2 ; F 2 <->A U E, F lt F 2 . The diagonal matrix elements exist (as we see from (2), Problem 1) for the states F x and F 2 . The sums A xx + sA yy + s*Azz, A xx + e?A yy + sA zz (e = e 2 "*/ 3 ) are transformed by the representation E. The selection rules for the non-diagonal elements are E<- >A lt A it E; Fi^-^Fi, F 2 ; F 2 <r^F 2 . The diagonal elements are non-zero for the states E, F lt F 2 . Problem 4. The same as Problem 3, but for symmetry D 3d . Solution. Azz is transformed by A ig , i.e. A Z z behaves as a scalar. The components A xx — A yy and A xy are transformed by E g ; the same is true of the components A X z, A yz . Decomposing the direct products of Eg with all the representations of the group Dad, we find the selection rules for the non-diagonal matrix elements ; E g <— > Aig, A 2g , E g ; E u <— > A 1U , A 2U , E u . The diagonal elements are not zero (as we see from (2), Problem 2) only for the states Eg and E u . 364 The Theory of Symmetry §98 §98. Continuous groups As well as the point groups enumerated in §93, there exist also what are called continuous point groups, having an infinite number of elements. These are the groups of axial and spherical symmetry. The simplest axial symmetry group is the group C m , which contains rota- tions C(<f>) through any angle <f> about the axis of symmetry ; this is called the two-dimensional rotation group. It may be regarded as the limiting case of the groups C n as n -> oo. Similarly, as limiting cases of the groups C nh , c nv> D n, D nh we obtain the continuous groups C mht C,,, D MI D xh . A molecule has axial symmetry only if it consists of atoms lying in a straight line. If it meets this condition, but is asymmetric about its midpoint, its point group will be the group C^, which, besides rotations about the axis, contains also reflections a v in any plane passing through the axis. If, on the other hand, the molecule is symmetrical about its midpoint, its point group will be D ooh = C mv xC t . The groups C ro , C wft , D^ cannot appear as the symmetry groups of a molecule. The group of complete spherical symmetry contains rotations through any angle about any axis passing through the centre, and reflections in any plane passing through the centre; this group, which we shall denote by K h , is the symmetry group of a single atom. It contains as a sub-group the group K of all spatial rotations (called the three-dimensional rotation group, or simply the rotation group). The group K h can be obtained from the group K by adding a centre of symmetry (K h — K x C f ). The elements of a continuous group may be distinguished by one or more parameters which take a continuous range of values. Thus, in the rotation group, the parameters might be the two angles determining the direction of the axis, and the angle of rotation about this axis. The general properties of finite groups described in §92, and the concepts appertaining to them (sub-groups, conjugate elements, classes, etc.), can be at once generalised to continuous groups. Of course, the statements which directly concern the order of the group (for instance, that the order of a sub- group divides the order of the group) are no longer meaningful. In the group C oov all planes of symmetry are equivalent, so that all reflec- tions a v form a single class with a continuous series of elements ; the axis of symmetry is bilateral, so that there is a continuous series of classes, each containing two elements C(±<£). The classes of the group D ooh are obtained at once from those of the group C mVi since D ooh = C aav xC i . In the rotation group K, all axes are equivalent and bilateral; hence the classes of this group are rotations through an angle of fixed absolute magnitude \<f>\ about any axis. The classes of the group K h are obtained at once from those of the group K. We have already found, in essence, the irreducible representations of the three-dimensional rotation group (without using the terminology of group theory), when determining the eigenvalues and eigenfunctions of the total §98 Continuous groups 365 angular momentum. For the angular momentum operator is (apart from a constant factor) the operator of an infinitely small rotation, and its eigenvalues characterise the behaviour of the wave functions with respect to spatial rotations. To a valued of the angular momentum there correspond 2/ + 1 different eigenfunctions ip] m > differing in the values of the component m of the angular momentum and all belonging to one (2/ + l)-fold degenerate energy level. Under rotations of the co-ordinate system, these functions are transformed into linear combinations of themselves, and thus give irreducible representations of the rotation group. Thus, from the group-theory point of view, the numbers j number the irreducible representations of the rotation group, and one (2/+ 1)- dimensional representation corresponds to each j. The number j takes integral and half- integral values, so that the dimension 2/+ 1 of the representation takes all the integral values 1, 2, 3, . . . . The base functions of these representations have been, in essence, investi- gated in Chapter VIII. The basis of a representation of given/ is formed by the 2j+ 1 independent components of a symmetrical spinor of rank 2/ (which are equivalent to the set of 2j+ 1 functions fym). The irreducible representations of the rotation group which correspond to half-integral values of j are distinguished by an important property. Under a rotation through 2tt, the base functions of the representations change sign (being components of a spinor of odd rank). Since, however, a rotation through 2n is the same as the unit element of the group, we reach the result that representations with half-integral j are, as we say, two-valued; to each element of the group (a rotation through an angle cf>, ^ ^ 2tt, about some axis) there correspond in such a representation not one but two matrices, with characters differing in sign.f An isolated atom has, as we have already remarked, the symmetry K h = KxC t . Hence, from the group-theory point of view, there corresponds to each term of the atom some irreducible representation of the rotation group K (determining the value of the total angular momentum J of the atom) and an irreducible representation of the group C^ (determining the parity of the state). J When the atom is placed in an external electric field, its energy levels are split. The number of different levels resulting and the symmetry of the corresponding states can be determined by the method described in §96. t It must be mentioned that "two-valued representations" of a group are not representations in the true sense of the word, since they are not given by one-valued base functions; see also §99. J Moreover, the Hamiltonian of the atom is invariant with respect to interchanges of the electrons. In the non-relativistic approximation, the co-ordinate and spin wave functions are separable, and we can speak of representations of the permutation group that are given by the co-ordinate functions. If the irreducible representation of the permutation group is given, the total spin S of the atom is deter- mined (§63). When the relativistic interactions are taken into account, however, the separation of the wave functions into co-ordinate and spin parts is not possible. The symmetry with respect to simul- taneous interchange of the co-ordinates and spins of the particles does not characterise the term, since Pauli's principle admits only those total wave functions which are antisymmetric with respect to all the electrons. This is in accordance with the fact that, when the relativistic interactions are taken into account, the spin is not, strictly speaking, conserved; only the total angular momentum / is conserved. 366 The Theory of Symmetry §98 It is necessary to decompose the (27+l)-dimensional representation of the symmetry group of the external field (given by the functions ifj JM ) into irreducible representations of this group. This requires a knowledge of the characters of the representation given by the functions iJjjm- By a rotation through an angle <£ about an axis the wave functions iftj M are, as we know, multiplied by e iM $, where M is the component of the angular momentum along that axis. The transformation matrix for the functions i/jjm will therefore be diagonal, with character ^4 e U J+l)$ _ e -u<t> > lJ) (<f>) = y e iM <t> = m=-J e W—\ or sin(7+A)(£ X {J) (<1>) = . J . (98.1) sinf<p With respect to inversion I, all the functions iff JM with different M behave in the same way, being multiplied by + 1 or - 1 according as the state of the atom is even or odd. Hence the character X {J) (I)= ±(27+1). (98.2) Finally, the characters corresponding to reflection in a plane a and rotary reflection through an angle <f> are found by writing these symmetry trans- formations as CT = /C 2 , S((f>)= IC(7T + <f>). Let us pause to consider also the irreducible representations of the axial symmetry group C mv . This problem has, in essence, been solved when we ascertained the classification of the electron terms of a diatomic molecule having this symmetry C aav (i.e. when the two atoms are different). To the terms 0+ and 0~ (with Q. = 0) there correspond two one-dimensional irreducible representations : the unit representation A\ and the representation A2, in which the base function is invariant under all rotations and changes sign under reflections in planes a Vy while to the doubly degenerate terms with Q, = 1, 2, ... there correspond two-dimensional representations denoted by Ei, E2, .... Under a rotation through an angle <f> about the axis, the base functions are multiplied by e ±iQ ^, while on reflection in planes a v they change into each other. The irreducible representations of the group D^ h = C oav xC i are ob- tained at once from those of the group C oav (and correspond to the classifi- cation of the terms of a diatomic molecule composed of like nuclei). If we take half-integral values for Q., the functions e ±in ^ give two-valued §99 Two-valued representations of finite point groups 367 irreducible representations of the group C^ v , corresponding to the terms of the molecule having half-integral spin.f §99. Two-valued representations of finite point groups To the states of a system with half-integral spin (and therefore half- integral total angular momentum) there correspond two-valued represen- tations of the point symmetry group of the system. This is a general property of spinors, and therefore holds for both continuous and finite point groups. The necessity thus arises of finding the two-valued irreducible representa- tions of finite point groups. As we have already remarked, the two-valued representations are not really true representations of a group. In particular, the relations discussed in §94 do not apply to them, and where all irreducible representations were considered in these relations (for example, in the relation (94.15) for the sum of the squared dimensions of the irreducible representations), only the true one-valued representations were meant. To find the two- valued representations, it is convenient to employ the following artifice (H. Bethe 1929). We introduce, in a purely formal manner, the concept of a new element of the group (denoted by Q) ; this is a rotation through an angle of 2tt about an arbitrary axis, and is not the unit element, but gives the latter when applied twice : Q 2 = E. Accordingly, rotations C n about the axes of symmetry of the «th order will give identical transformations only after being applied 2w times (and not n times) : C n n = Q, C n *" = E. (99.1) The inversion /, being an element which commutes with all rotations, must give E as before on being applied twice. A twofold reflection in a plane, however, gives Q, not E: g 2 = Q, c* = E; (99.2) this follows, since the reflection can be written in the form a h = IC % . As a result we obtain a set of elements forming some fictitious point symmetry group, whose order is twice that of the original group ; such groups we shall call double point groups. The two-valued representations of the actual point group will clearly be one-valued (i.e. true) representations of the correspond- ing double group, so that they can be found by the usual methods. The number of classes in the double group is greater than in the original group (but not, in general, twice as great). The element Q commutes with all f Contrary to the result for the three-dimensional rotation group, it would here be possible, by a suitable choice of fractional values of CI, to obtain not only one- valued and two-valued representations, but also those of three or more values. However, the physically possible eigenvalues of the angular momentum, which is the operator of an infinitely small rotation, are determined by the representations of the aforementioned three-dimensional rotational group. Hence the three (or more)-valued repre- sentations of the two-dimensional rotation group (and of any finite symmetry group), though mathe- matically determinate, are without physical significance. 368 The Theory of Symmetry §99 the other elements of the group,f and hence always forms a class by itself. If the axis of symmetry is bilateral, the elements C n k and C n 2n ~ k = QC n n ~ k are conjugate in the double group. Hence, when axes of the second order are present, the distribution of the elements among classes depends also on whether these axes are bilateral (in ordinary point groups this is unimportant, since C 2 is the same as the opposite rotation C 2 -1 ). Thus, for instance, in the group T the axes of the second order are equiva- lent, and each of them is bilateral, while the axes of the third order are equiva- lent but not bilateral. Hence the 24 elements of the double group J T' are distributed in seven classes : E, Q, the class of three rotations C 2 and three C 2 Q, and the classes 4C 3 , 4C 3 2 , 4C 3 £, 4C 3 2 £. The irreducible representations of a double point group include, firstly, representations which are the same as the one- valued representations of the simple group (a unit matrix corresponding to both Q and E)\ secondly, the two-valued representations of the simple group, a negative unit matrix corresponding to Q. It is these latter representations in which we are now interested. The double groups C n ' (n = 1, 2, 3, 4, 6) and 5 4 ', like the corresponding simple groups, are cyclic. || All their irreducible representations are one- dimensional, and can be found without difficulty as shown in §95. The irreducible representations of the groups D n ' (or C nv ', which are isomorphous with them) can be found by the same method as for the cor- responding simple groups. These representations are given by functions of the form e ±ik ^ y where <j> is the angle of rotation about an axis of the nth order, and k is given half-integral values (the integral values correspond to the ordinary one-valued representations). Rotations about horizontal axes of the second order change these functions into one another, while the rota- tion C n multiplies them by e ±2 " ik / n . It is a little less easy to find the representations of the double cubic groups. The 24 elements of the group T" are divided among seven classes. Hence there are altogether seven irreducible representations, of which four are the same as those of the simple group T. The sum of the squared dimensions of the remaining three representations must be 12, and hence we find that they are all two-dimensional. Since the elements C 2 and C 2 Q belong to the same class, x(C 2 ) = xiQQ) = ~ x(Q)> whence we conclude that x(Q) = in all three representations. Next, at least one of the three representations must be real, since complex representations can occur only in conjugate pairs. Let us consider this representation, and suppose that the matrix of the element C 3 is brought to diagonal form, with diagonal elements a 1} a^. Since C 3 3 = Q, a x 3 = a 2 3 = — 1. In order that x(Cs) — a i+ a z ma y De rea l> we must take a x = e ni / 3 , a 2 = r^/ 3 . Hence we find that x(C 3 ) = 1, x(Q a ) "f This is obvious for rotations and inversion; for a reflection in a plane, it follows since the reflec- tion can be represented as the product of an inversion and a rotation. J We distinguish the double groups by primes to the symbols for the ordinary groups. || The groups S 2 ' = C/, S 6 ' = C 8 ,', however, which contain the inversion J, are Abelian but not cyclic. §99 Two-valued representations of finite point groups 369 = a ^-\-a£ = — 1. Thus one of the required representations is obtained. By comparing its direct products with the two complex conjugate one-dimensional representations of the group T, we find the other two representations. By means of similar arguments, which we shall not pause to give here, we may find the representations of the group O'. Table 8 gives the characters of the representations of the double groups mentioned above. Only those representations are shown which correspond to two-valued representations of the ordinary groups. The isomorphous double groups have the same representations. The remaining point groups are isomorphous with those we have con- sidered, or else are obtained by direct multiplication of the latter by the group C t , so that their representations do not need to be specially calculated. For the same reasons as for ordinary representations, two complex con- jugate two-valued representations must be regarded, from the physical point of view, as one representation of twice the dimension. It is necessary to pair one-dimensional two-valued representations even when they have real characters. For (see §60) in systems with half-integral spin, complex conjugate wave functions are linearly independent. Hence, if we have a two- valued one- dimensional representationf with real characters (given by some function if/), then, although the complex conjugate function i/j* is transformed by the same representation, we can nevertheless see that ifj and ^r* are linearly independent. Since, on the other hand, the complex conjugate wave functions must belong to the same energy level, we see that in physical applications this representation must be doubled. PROBLEM Determine how the levels of an atom (with given values of the total angular momentum J) are split when it is placed in a field having the cubic symmetry J O. Solution. The wave functions of the states of an atom with angular momentum / and various values Mj give a (2/+l)-dimensional reducible representation of the group O, with characters determined by the formula (98.1). Decomposing this representation into irreduc- ible parts (one-valued for integral / and two-valued for half-integral /), we at once find the required splitting (cf. §96). We shall list the irreducible parts of the representations corres- ponding to the first few values of J: 7=0 A 1 1/2 E{ 1 *"i 3/2 G' 2 E+F 2 5/2 E 2 '+G' 3 A 2 +F 1 +F 2 t Such representations are found in the group C n ' for odd n; the characters are x(C„*) = (—!)*• % For example, an atom in a crystal lattice. The presence or absence of a centre of symmetry in the symmetry group of the external field is immaterial to this problem, since the behaviour of the wave function on inversion (the parity of the level) is unrelated to the angular momentum J. 370 The Theory of Symmetry §99 Table 8 Two-valued representations of point groups A' E Q C 2 <*> C 2 <*>0 C 2 <*> C 2 <">£> c 2 <*> C 2 < 2 >£> E' 2 -2 *>»' E Q c 3 c 3 2 c 3 p 3C/ 2 3C/ 2 .,| 1 1 -l -l -1 -1 1 1 i — i —i i E % ' 2 -2 1 -1 D* E c 2 C 2 Q c 3 CJQ c 3 2 C 3 Q c 6 C 6 6 c 6 5 C 6 £ 3C/ 2 3C7 2 3C7' 2 3C7' 2 Ex' 2 -2 1 -1 V3 -V3 E % ' 2 -2 1 -1 -V3 V3 E s ' 2 -2 -2 2 d; E c 2 C 2 c 4 C7Q C 4 3 C 4 Q 2£/ 2 2TJ\ 2U\Q E x ' 2 -2 V2 -V2 E 2 ' 2 -2 -V2 V2 T E 4C, 4C 3 2 4C 3 Q 4C 3 2 3C 2 3C 2 Q E' 2 -2 1 -1 -1 1 °'l 2 -2 e -e 2 — s e 2 { 2 -2 e 2 — s -e 2 s O' E Q 4C 3 4C 3 *£> 4C 3 2 4C 3 £> 3C 4 2 3C 4 2 3C 4 3C 4 3 3C 4 3 3C 4 6C 2 ec 2 Q E^ 2 -2 1 -1 V2 -V2 E t ' 2 -2 1 -1 -V2 V2 G' 4 -4 -1 1 CHAPTER XIII POLYATOMIC MOLECULES §100. The classification of molecular vibrations In its applications to polyatomic molecules, group theory first of all resolves at once the problem of the classification of their electron terms, i.e. of the energy levels for a given situation of the nuclei. They are classified according to the irreducible representations of the point symmetry group appropriate to the configuration of the nuclei. Here, however, we must emphasise what is really obvious, that the classification thus obtained belongs to the definite nuclear configuration considered, since the symmetry is in general destroyed when the nuclei are displaced. We usually discuss the configuration cor- responding to the equilibrium position of the nuclei. In this case the classi- fication continues to possess a certain amount of meaning even when the nuclei execute small vibrations, but of course becomes meaningless when the vibrations can no longer be regarded as small. In the diatomic molecule this question did not arise, since its axial sym- metry is of course preserved under any displacement of the nuclei. A similar situation occurs for triatomic molecules also. The three nuclei always lie in a plane, which is a plane of symmetry of the molecule. Hence the classification of the electron terms of the triatomic molecule with respect to this plane (wave functions symmetric or antisymmetric with respect to reflection in the plane) is always possible. For the normal electron terms of polyatomic molecules there is an empirical rule according to which, in the great majority of molecules, the wave function of the normal electron state is completely symmetrical (this rule, for diatomic molecules, has already been mentioned in §78). In other words, the wave function is invariant with respect to all the elements of the symmetry group of the molecule, i.e. it belongs to the unit irreducible representation of the group. The application of the methods of group theory is particularly significant in the investigation of molecular vibrations (E. Wigner 1930). Before beginning a quantum-mechanical investigation of this problem, a purely classical discussion of the vibrations of the molecule is necessary, in which it is regarded as a system of several interacting particles (the nuclei). A system of N particles (not lying in a straight line) has 3N— 6 vibrational degrees of freedom; of the total number of degrees of freedom 3N, three correspond to translational and three to rotational motion of the system as a wholef. The energy of a system of particles executing small vibrations t See Mechanics, §§23, 24. If all the particles lie in a straight line, the number of vibrational degrees of freedom is 3N— 5 ; in this case, only two co-ordinates correspond to rotation, since it is meaningless to speak of the rotation of a linear molecule about its axis. 13 371 372 Polyatomic Molecules §100 can be written E = * Be "^"fc+i 5 kikUiUk > (100. 1) where m ik , k ik are constant coefficients, and the u t are the components of the vector displacements of the particles from their equilibrium positions (the suffixes *, k denumerate both the components of the vector and the particles). By a suitable linear transformation of the quantities u t , we can eliminate from (100.1) the co-ordinates corresponding to translational motion and rotation of the system, and take the vibrational co-ordinates in such a way that both the quadratic forms in (100.1) are transformed into sums of squares. Normalising these co-ordinates so as to make all the coefficients in the expression for the kinetic energy unity, we obtain the vibrational energy in the form E = i s QJ+i S <o a 2 f QJ. (100.2) The vibrational co-ordinates Q ai are said to be normal; the co a are the fre- quencies of the corresponding independent vibrations. It may happen that the same frequency (which is then said to be multiple) corresponds to several normal co-ordinates ; the suffix a to the normal co-ordinate gives the number of the frequency, and the suffix i = 1, 2, ... ,/ a numbers the co-ordinates belonging to a given frequency (J^ being the multiplicity of the frequency). The expression (100.2) for the energy of the molecule must be invariant with respect to symmetry transformations. This means that, under any transformation belonging to the point symmetry group of the molecule, the normal co-ordinates Q ai , i = 1, 2, ... ,f a (for any given a) are transformed into linear combinations of themselves, in such a way that the sum of the squares £ Q ai 2 remains unchanged. In other words, the normal co-ordinates i belonging to any particular eigenfrequency of the vibrations of the molecule give some irreducible representation of its symmetry group ; the multiplicity of the frequency determines the dimension of the representation. The irreducibility follows from the same considerations as were given in §96 for the solutions of Schrodinger's equation. The equality of the frequencies corresponding to two different irreducible representations would be an improbable coincidence. An exception is again formed by the irreducible representations with complex conjugate systems of characters. Since the normal co-ordinates are by their physical nature real quantities, two complex conjugate representations correspond physically to one eigenfrequency of twice the multiplicity. These considerations enable us to carry out a classification of the eigen- vibrations of a molecule without solving the complex problem of actually determining its normal co-ordinates. To do so, we must first find (by the method described below) the representation given by all the vibrational co-ordinates together, which we shall call the total representation; this §100 The classification of molecular vibrations 373 representation is reducible, and on decomposing it into irreducible parts we determine the multiplicities of the eigenfrequencies and the symmetry properties of the corresponding vibrations. Here it may happen that the same irreducible representation appears several times in the total represen- tation; this means that there are several different frequencies of the same multiplicity and with vibrations of the same symmetry. To find the total representation, we start from the fact that the characters of a representation are invariant with respect to a linear transformation of the base functions. Hence they can be calculated by using as base functions not the normal co-ordinates, but simply the components u t of the vectors of the displacements of the nuclei from their equilibrium positions. First of all, it is evident that, to calculate the character of some element G of a point group, we need consider only those nuclei which (or, more exactly, whose equilibrium positions) remain fixed under the given symmetry trans- formation. For if, under the rotation or reflection G in question, nucleus 1 is moved to a new position, previously occupied by a similar nucleus 2, this means that under the operation G a displacement of nucleus 1 is trans- formed into a displacement of nucleus 2. In other words, there will be no diagonal elements in the rows of the matrix G ik which correspond to this nucleus (i.e. to its displacement u { ). The components of the displacement vector of a nucleus whose equilibrium position is not affected by the operation G, on the other hand, are evidently transformed into combinations of them- selves, so that they may be considered independently of the displacement vectors of the remaining nuclei. Let us first consider a rotation C{<j>) through an angle <£ about some sym- metry axis. Let u x , u y , u z be the components of the displacement vector of some nucleus, whose equilibrium position is on the axis, and hence is unaffected by the rotation. Under the rotation these components are transformed, like those of any ordinary (polar) vector, according to the for- mulae (the #-axis being the axis of symmetry) u 'x = u x cos<£+#„ sin<£, u' y = —u x sm<f>-\-u v cos(f> t u\ =11,. The character, i.e. the sum of the diagonal terms of the transformation matrix, is 1+2 cos<f>. If altogether N c nuclei lie on the axis in question, the total character is N c (l+2cos<f>). (100.3) However, this character corresponds to the transformation of all the 3iV displacements u t ; hence it is necessary to separate the part corresponding to the transformations of translation and (small) rotation of the molecule as a whole. The translation is determined by the displacement vector U of the centre of mass of the molecule ; the corresponding part of the character is 374 Polyatomic Molecules §100 therefore 1+2 cos^. The rotation of the molecule as a whole is determined by the vector 8SI of the angle of rotation.f The vector 8£l is axial, but with respect to rotations of the co-ordinate system an axial vector behaves like a polar vector. Hence a character of 1 +2 cos <j> also corresponds to the vector 8€l. Altogether, therefore, we must subtract from (100.3) a quantity 2(1 +2 cos<£). Thus we finally have the character x(C) of the rotation C(<f>) in the total vibrational representation : X (C) = (7Vc-2)(l+2 cos<£). (100.4) The character of the unit element is evidently just the total number of vibrational degrees of freedom: x( E ) = 3JV-6 (as is obtained from (100.4) when N c = N, <f> = 0). In an exactly similar manner, we calculate the character of the rotary- reflection transformation S(<f>) (a rotation through an angle <f> about the ^-axis and a reflection in the xy-plane). Here a vector is transformed accord- ing to the formulae u' x = u x cos0+Mj, sin<£, u' y = — u x sm(f>+Uy COS(f>, u'z = — u z, to which there corresponds a character — 1+2 cos <f>. Hence the character of the representation given by all the 3AT displacements u i is iV,s(-l+2cos<£), (100.5) where N s is the number of nuclei left unmoved by the operation S(<j>) ; this number is evidently either none or one. To the vector U of the displacement of the centre of mass there corresponds a character —1+2 cos<f>. The vector 8SI, being an axial vector, is unchanged by an inversion of the co-ordinate system ; on the other hand, the rotary-reflection transformation S(<f>) can be represented in the form S(cf>) = C(<f>)a h = C{4>)C 2 I = C(tt+0)/, i.e. as a rotation through an angle 7T+0, followed by an inversion. Hence the character of the transformation S(<f>) applied to the vector 8SI is equal to the character of the transformation C{tt-\-(J>) applied to an ordinary vector, i.e. it is 1 +2 cos (it +<£) = 1 — 2 cos <f>. The sum ( — 1 +2 cos <f>) +(1 —2 cos cf>) = 0, so that we reach the conclusion that the expression (100.5) is equal to the required character x(S) of the rotary-reflection transformation S((f>) in the total representation : x (5) = Ns(-l+2 cos<£). (100.6) In particular, the character of reflection in a plane (<£ = 0) is x( CT ) = N ai while that of an inversion (<f> = rr) is xCO = — 3AT 7 . f As is well known, the angle of a small rotation can be regarded as a vector 8SI, whose modulus is equal to the angle of rotation and which is directed along the axis of rotation in the direction deter- mined by the corkscrew rule. The vector 8£l so denned is clearly axial. §100 The classification of molecular vibrations 375 Having thus determined the characters x °f tne total representation, we have only to decompose it into irreducible representations, which is done by formula (94.14) and the character tables given in §95 (see the Problems at the end of the present section). To classify the vibrations of a linear molecule there is no need to have recourse to group theory. The total number of vibrational degrees of freedom is 3iV— 5. Among the vibrations, we must distinguish those in which the atoms remain in a straight line, and those where this does not happen, f The number of degrees of freedom in the motion of N particles in a straight line is N; of these, one corresponds to the translational motion of the molecule as a whole. Hence the number of normal co-ordinates of the vibrations which leave the atoms in a straight line is N— 1; in general, N— 1 different eigenfrequencies correspond to them. The remaining (3iV — 5)— (N— 1) = 22V — 4 normal co-ordinates relate to vibrations which destroy the col- linearity of the molecule; to these, there correspond N—2 different double frequencies (two normal co-ordinates, corresponding to the same vibrations in two mutually perpendicular planes, belong to each frequency). J PROBLEMS Problem 1. Classify the normal vibrations of the molecule NH 3 (an equilateral triangular pyramid, with the N atom at the vertex and the H atoms at the corners of the base; Fig. 41). Solution. The point symmetry group of the molecule is C Sv . Rotations about an axis of the third order leave only one atom (N) fixed, while reflections in planes each leave two atoms fixed (N and one H). From formulae (100.4), (100.6) we find the characters of the total representation : E lCo 3 g„ 6 2 Decomposing this representation into irreducible parts, we find that it contains the repre- sentations A x and E twice each. Thus there are two simple frequencies corresponding to vibrations of the type A u which conserve the complete symmetry of the molecule (what are called totally symmetric vibrations), and two double frequencies with corresponding normal co-ordinates which are transformed into combinations of each other by the representation E. Problem 2. The same as Problem 1, but for the molecule H 2 (Fig. 42). Solution. The symmetry group is C 2v . The transformation C 2 leaves the O atom fixed ; the transformation o„ (a reflection in the plane of the molecule) leaves all three atoms fixed ; f If the molecule is symmetrical about its centre, a further characteristic of the vibrations appears ; see Problem 10 at the end of this section. J Using the notation for the irreducible representations of the group C KV (see §98), we can say that there are N— 1 vibrations of the type A lt and N — 2 of the type £^. 376 Polyatomic Molecules §100 Fig. 42 the reflection a' v leaves only the O atom fixed. The characters of the total representation are E Co a„ a'., 3 13 1 This representation divides into the irreducible representations 2A X , \B lt i.e. there are two totally symmetric vibrations and one with the symmetry given by the representation B 1 ; all the frequencies are simple. Fig. 42 shows the corresponding normal vibrations. Problem 3. The same as Problem 1, but for the molecule CHC1 3 (Fig. 43a). Solution. The symmetry group of the molecule is C 3v . By the same method we find that there are three totally symmetric vibrations A x and three double vibrations of the type E. Problem 4. The same as Problem 1 , but for the molecule CH 4 (the C atom is at the centre of a tetrahedron with the H atoms at the vertices; Fig. 43b). Solution. The symmetry of the molecule is T d . The vibrations are 1A U IE, 2F 2 . Problem 5. The same as Problem 1, but for the molecule C 6 H 6 (Fig. 43c). Solution. The symmetry of the molecule is D th . The vibrations are 2A ig , lA zg , lA iU , lB ig , lBiu, \B %g , 3B2U, lExg, oEiu, ^E^g, 2E 2 U- Problem 6. The same as Problem 1, but for the molecule OsF 8 (the Os atom is at the centre of a cube with the F atoms at the vertices; Fig. 43d). Solution. The symmetry of the molecule is O^. The vibrations are lA ig , XA^u, \Eg, 1E U , 2i ? 1 u> 2F 2 g, 1jF2U. Problem 7. The same as Problem 1 , but for the molecule UF « (the U atom is at the centre of an octahedron with the F atoms at the vertices; Fig. 43e). Solution. The symmetry of the molecule is Of,. The vibrations are lA ig , \E g , 2F 1U , §100 The classification of molecular vibrations » H H 377 (a) (d) y y F. Os o V y A (g) Fig. 43 Problem 8. The same as Problem 1, but for the molecule QHg (Fig. 43 f). Solution. The symmetry of the molecule is D M . The vibrations are ZA xg , 1A 1U , 2A iU , "SEg, 2Eu- Problem 9. The same as Problem 1, but for the molecule C 2 H 4 (Fig. 43g; all the atoms are coplanar). Solution. The symmetry of the molecule is D 2h . The vibrations are 3A ig , IA 1U , 2B ig , \B 1U , 2B au , 1B 2 g, 2B 2U ', the axes of co-ordinates are taken as shown in the figure. Problem 10. The same as Problem 1, but for a linear molecule of N atoms symmetrical about its centre. Solution. To the classification of the vibrations of a linear molecule considered in the text, we must add the classification from the behaviour with respect to inversion in the centre. There are two distinct cases, according as JV is even or odd. If N is even (JV = 2p), there is no atom at the centre of the molecule. On giving to the p atoms in one half of the molecule independent displacements along the line, and to the re- maining p atoms equal and opposite displacements, we find that p of the vibrations leaving the atoms in line are symmetrical with respect to the centre, while the remaining (2p — 1) — p = p—\ vibrations of this type are antisymmetrical. Next, p atoms have 2p degrees of freedom for motions in which the atoms do not remain in line. On giving equal and opposite displace- ments to symmetrically placed atoms, we should obtain 2p symmetrical vibrations; of these, however, the two corresponding to a rotation of the molecule must be removed. Thus there are p — 1 double frequencies of vibrations which bring the atoms out of line and are symmetri- cal about the centre, and the same number [(2p— 2)— (p— 1) = p— 1] which are antisym- metrical. Using the notation for the irreducible representations of the group D<x> h (see 378 Polyatomic Molecules §101 the end of §98), we can say that there are p vibrations of the type A ig and p—\ of the types ■A-xu, E\g, Em- it N is odd (N = 2p + l), similar arguments show that there are p vibrations of each of the types A ig , A 1U , E 1U and p — 1 of the type E ig . §101. Vibrational energy levels From the viewpoint of quantum mechanics, the vibrational energy of a molecule is determined by the eigenvalues of the Hamiltonian f f &* = | S .2 4< 2 +i J aJ&QJ, (101.1) where P^i = —^I^Qoci are tne momentum operators corresponding to the normal co-ordinates Q ai . Since this Hamiltonian falls into the sum of in- dependent terms K^t^+^Ga* 2 )* tne ener gy levels are given by the sums ES* = h S co a S K,+i) = S /KK+iU (101.2) where ^ a = S ^ ai , and / a is the multiplicity of the frequency a> a . The wave functions are given by the products of the corresponding wave functions for linear harmonic oscillators : 0=n0„ (101.3) a where «/r a = constant xexp{-|c a 2 S ^ ai 2 }n H v Jc„<Qaa)* (101.4) where i/„ denotes the Hermite polynomial of order v, and c a = ^(wjh). If there are multiple frequencies among the a> a , the vibrational energy levels are in general degenerate. The energy (101.2) depends only on the sums v a = S v ai . Hence the degree of degeneracy of the level is equal to the number of ways of forming the given set of numbers v a from the v a{ . For a single number v a it isj- (f«+/«-i)W(/«-i)i. Hence the total degree of degeneracy is n — (101.5) >.!(/«-l)l For double frequencies, the factors in this product are v x + l, while for triple frequencies they are i(^ a +l)(^ a +2). It must be borne in mind that this degeneracy occurs only so long as we consider purely harmonic vibrations. When terms of higher order in the normal co-ordinates are taken into account in the Hamiltonian {anharmonic vibrations), the degeneracy is in general removed, though not completely (see §104 for a further discussion of this point). t This is the number of ways in which » a balls can be distributed among / a urns. §101 Vibrational energy levels 379 The wave functions (101.3) belonging to the same degenerate vibrational term give some representation (in general reducible) of the symmetry group of the molecule. The functions belonging to different frequencies are transformed independently of one another. Hence the representation given by all the functions (101.3) is the product of the representations given by the functions (101.4), so that we need consider only the latter. The exponential factor in (101.4) is invariant with respect to all the symmetry transformations. In the Hermite polynomials, the terms of any given degree are transformed only into similar terms; a symmetry trans- formation evidently does not change the degree of any term. Since, on the other hand, each Hermite polynomial is completely determined by its highest term, it follows that it is sufficient to consider only the highest term, writing /« T^HiJc&rt) = constant x& a M>e* r «' - Qocf /«'«+ + terms of lower degree. The functions for which the sum v a = 2 v ai has the same value belong to the same term. Thus we have a representation given by the products of v a quantities Q ai ; this is just the symmetric product (see §94) of the irreducible representation given by the Q ai with itself v a times. For one-dimensional representations, the finding of the characters of their symmetric products with themselves v times is trivial :f Xv (G) = [x(G)]». For two- and three-dimensional representations it is convenient to use the following mathematical device. J The sum of the squared base functions of an irreducible representation is invariant with respect to all symmetry transformations. Hence we can formally regard these functions as the com- ponents of a vector in two or three dimensions, and the symmetry transfor- mations as some rotations (or reflections) applied to these vectors. We emphasise that there is in general no relation between these rotations and reflections and the actual symmetry transformations, the former depending (for any given element G of the group) also on the particular representation considered. Let us consider two-dimensional representations more closely. Let x(^) be the character of some element of the group in the two-dimensional repre- sentation concerned, with x(G) # 0. The sum of the diagonal elements of the transformation matrix for the components x, y of a two-dimensional vector on rotation through an angle <j> in a plane is 2 cos <f>. Putting 2cos<f> = X (G), (101.6) we find the angle of the rotation which formally corresponds to the element f We use the notation X v (&) * n place of the cumbersome [x c ] (G). X It was applied to this problem by A. S. Kompaneets (1940). 380 Polyatomic Molecules §102 G in the irreducible representation considered. The symmetric product of the representation with itself v times is the representation whose basis is formed by the v + 1 quantities *", x^^y, ... , y v . The characters of this representation aref Xv (G) =sin(*>+l)#5in0. (101.7) The case where x(G) = requires special consideration, since a zero charac- ter corresponds both to a rotation through \n and to a reflection. If x(G 2 ) = —2, we have a rotation through \tt, and for Xv(G) we obtain X*(G) = — *[l + (— 1)']. (101.8) If x(G 2 ) — 2, on the other hand, x(G) must be regarded as the character of a reflection (i.e. a transformation x -> x, y -> — y) ; then xv(G)=m+(-m- (ioi.9) We can similarly obtain the formulae for the symmetric products of three- dimensional representations. The finding of the rotation or reflection which formally corresponds to an element of the group in a given representation is easily accomplished with the aid of Table 7. This is the transformation which corresponds to the given x(G) in that isomorphous group in which the co-ordinates are transformed by the representation in question. Thus, for the representation F t of the groups O and T d we must take a transformation from the group O, but for the representation F 2 we must take one from the group T d . We shall not pause here to derive the corresponding formulae for the characters Xv(G)- §102. Stability of symmetrical configurations of the molecule For a symmetrical position of the nuclei, an electron term of the molecule may be degenerate, if there are among the irreducible representations of the symmetry group one or more whose dimensions exceed unity. We may ask whether such a symmetrical configuration is a stable equilibrium configura- tion of the molecule. Here we shall entirely neglect the effect of spin (if any), since this effect is usually insignificant in polyatomic molecules. The degene- racy of the electron terms of which we shall speak is therefore only the "orbi- tal" degeneracy, and is unrelated to the spin. If the configuration in question is stable, the energy of the molecule as a function of the distances between the nuclei must be a minimum for the given position of the nuclei. This means that the change in the energy due to a small displacement of the nuclei must contain no terms linear in the displacements. *|* For purposes of calculation it is convenient to take the base functions in the form (x+iy) v , {x+iyY'^x—iy), ... , (x— iy) v ; the matrix of the rotation is then diagonal, and the sum of the diagonal elements takes the form §102 Stability of symmetrical configurations of the molecule 381 Let fi be the Hamiltonian of the electron state of the molecule, the distances between the nuclei being regarded as parameters. We denote by i? the value of this Hamiltonian for the symmetrical configuration considered. The quantities defining the small displacements of the nuclei can be taken as the normal vibrational co-ordinates Q^. The expansion of A in powers of the Q ai is of the form tf = # + S V a & ai +J i W 0liM Q 0li Q fik + ... . (102.1) The expansion coefficients V, W, ... are functions only of the co-ordinates of the electrons. Under a symmetry transformation, the quantities Q ai are transformed into combinations of one another, and the sums in (102.1) are changed into other sums of the same form. Hence we can formally regard the symmetry transformation as a transformation of the coefficients in these sums, the Q ai remaining unchanged. Here, in particular, the coefficients V ai (for any given a) will be transformed by the same representation of the symmetry group as the corresponding co-ordinates Q ai . This follows at once from the fact that, by virtue of the invariance of the Hamiltonian under all symmetry transformations, the group of terms of any given order in its expansion must be invariant also, and in particular the linear terms must be invariant.! Let us consider some electron term E Q which is degenerate in the sym- metrical configuration. A displacement of the nuclei which destroys the symmetry of the molecule generally results in a splitting of the term. The amount of the splitting is determined, as far as terms of the first order in the displacements of the nuclei, by the secular equation formed from the matrix elements of the linear term in the expansion (102.1), V, -g&iJlM^d* (102.2) where tp p , tp a are the wave functions of electron states belonging to the degenerate term in question (and are chosen to be real). The stability of the symmetrical configuration requires that the splitting linear in Q should be zero, i.e. all the roots of the secular equation must vanish identically. This means that the matrix V pa must itself be zero. Here, of course, we must consider only those normal vibrations which destroy the symmetry of the molecule, i.e. we must omit the totally symmetric vibrations (which corres- pond to the unit representation of the group). Since the Q ai are arbitrary, the matrix elements (102.2) vanish only if all the integrals t Strictly speaking, the quantities F" a ,- must be transformed by the representation which is the com- plex conjugate of that by which the Q a i are transformed. However, as we have already pointed out, if two complex conjugate representations are not the same, they must physically be considered to- gether as one representation of twice the dimension. The above remark is therefore unimportant. 382 Polyatomic Molecules §102 vanish. Let Z) el be the irreducible representation by which the electron wave functions ip p are transformed, and D a the same for the quantities V ai ; as we have already remarked, the representations D a are those by which the corresponding normal co-ordinates Q ai are transformed. According to the results of §97, the integrals (102.3) will be non-zero if the product [£)(ei)2j xD a contains the unit representation or, what is the same thing, if |7)(ei)2j contains D a . Otherwise all the integrals vanish. Thus a symmetrical configuration is stable if the representation [Z> (el > 2 ] does not contain any (except the unit representation) of the irreducible representations D a which characterise the vibrations of the molecule. This condition is always satisfied for non-degenerate electron states, since the sym- metric product of a one- dimensional representation with itself is the unit representation. Let us consider, for instance, a molecule of the type CH 4 , in which one atom (C) is at the centre of a tetrahedron, with four atoms (H) at the vertices. This configuration has the symmetry T d . The degenerate electron terms correspond to the representations E, F lt F 2 of this group. The molecule has one normal vibration A x (a totally symmetric vibration), one double vibration E, and two triple vibrations F 2 (see §100, Problem 4). The symmetric pro- ducts of the representations E, F^ F 2 with themselves are [£"] = A ± +E, [Ffl = [FJ] = A,+E+F 2 . We see that each of these contains at least one of the representations E, F 2 , and hence the tetrahedral configuration considered is unstable when there are degenerate electron states. This result constitutes a general rule (H. A. Jahn and E. Teller 1937). A more thorough investigation! than that given above, considering all possible types of symmetrical configuration of the nuclei, shows that, when there is a degenerate electron state, any symmetrical position of the nuclei (except when they are collinear) is unstable. As a result of this instability, the nuclei move in such a way that the symmetry of their configuration is destroyed, the degeneracy of the term being completely removed. In particular, we can say that the normal electron term of a symmetrical (non-linear) molecule can only be non-degenerate. As we have just mentioned, the linear molecules alone form an exception. This is easily seen, without using group theory. A displacement of a nucleus whereby it moves off the axis of the molecule is an ordinary vector with £ and r] components (the £-axis being along the axis of the molecule). We have seen in §87 that such vectors have matrix elements only for transitions in which the angular momentum A about the axis changes by unity. On the other hand, to a degenerate term of a linear molecule there correspond states with angular momenta A and —A about the axis (A ^ 1). A transition be- tween them changes the angular momentum by at least 2, and therefore the t See H. A. Jahn and E. Teller, Proceedings of the Royal Society A 161, 220, 1937. §103 Quantisation of the rotation of a rigid body 383 matrix elements always vanish. Thus the linear position of the nuclei in the molecule may be stable, even if the electron state is degenerate. §103. Quantisation of the rotation of a rigid body The investigation of the rotational levels of a polyatomic molecule is often hampered by the necessity of considering the rotation simultaneously with the vibrations. As a preliminary example, let us consider the rotation of a molecule as a solid body, i.e. with the atoms "rigidly fixed". Let £, r], I be a system of co-ordinates with axes along the three principal axes of 'inertia of a rigid body, and rotating with it. The corresponding Hamiltonian is obtained by replacing the components J ^ J v J^ of the angular momentum of the rotation, in the classical expression for the energy, by the corresponding operators : g =W (ll+ll + Jl\ (103.1) \ I A Ib 1C / where I A , I B , I c are the principal moments of inertia of the body. The commutation rules for the operators J^J V ,J C of the angular momen- tum components in a rotating system of co-ordinates are not obvious, since the usual derivation of the commutation rules relates to the components Jx> Jv> Jz in a fixed system of co-ordinates. They are, however, easily obtained by using the formula (J.a)(J.b)-(J.b)(J.a) = -/J.axb, (103.2) where a, b are any two commuting vectors which characterise the body in question. This formula is easily verified by calculating the left-hand side of the equation in the fixed system of co-ordinates *, y, z, using the general rules for the commutation of angular momentum components with one another and with the components of an arbitrary vector. Let a and b be unit vectors along the £ and y axes. Then axb is a umt vector along the £-axis, and (103.2) gives UJ,-UJi = -ft- < 103 - 3 > Two other relations are obtained similarly. Thus we reach the result that the commutation rules for the operators of the angular momentum com- ponents in the rotating system of co-ordinates differ from those in a fixed system only in the sign on the right-hand side of the equation. Hence it follows that all the results which we have previously obtained from the com- mutation rules, relating to the eigenvalues and matrix elements, hold for j.j y j also, with the difference that all expressions must be replaced by their complex conjugates. In particular, the eigenvalues of / ? (as of J z ) are the integers k = —J, ...,/• The finding of the eigenvalues of the energy of a rotating body (a top, 384 Polyatomic Molecules §103 as it is called) is simplest for the case where all three principal moments of inertia of the body are equal: I A = I B = I C = I ( a spherical top). This holds for a molecule in cases where it has the symmetry of one of the cubic point groups. The Hamiltonian (103.1) takes the form fi = frp/21, and its eigenvalues are E = &J(J+1)I2I. (103.4) Each of these energy levels is degenerate with respect to the 2J+ 1 directions of the angular momentum relative to the body itself (i.e. with respect to the values of J%= A).| There is also no difficulty in calculating the energy levels in the case where only two of the moments of inertia of the body are the same: I A — I B ^ I (a symmetrical top). This holds for molecules having one axis of symmetry of order above the second. The Hamiltonian (103.1) takes the form # = k 2 (j£ 2 +J v 2 )I2I A +H*Jfl2I c = Wl2I A+ &(L-l>)]t. (103 . 5) Hence we see that, in a state with given values of/ and k, the energy is E .^-JU+Vi+wQ-Ly, (103.6) which determines the energy levels of a symmetrical top. The degeneracy with respect to values of k which occurred for a spherical top is here partly removed. The values of the energy are the same only for values of k differing in sign alone, corresponding to opposite directions of the angular momentum relative to the axis of the top. Thus the energy levels of a symmetrical top are (for k # 0) doubly degenerate. For I A # I B # I c (an asymmetrical top), the calculation of the energy levels in a general form is impossible. The degeneracy with respect to the directions of the angular momentum relative to the body is here removed completely, so that 2J+ 1 different non- degenerate levels correspond to any given /. The calculation of these levels involves the solution of Schrod- inger's equation in matrix form; this amounts to solving a secular equation of degree 2J+ 1, formed from the matrix elements Hj\. with the given value of/ (O. Klein 1929). The matrix elements Hj% are defined with respect to the wave functions ip Jk of states in which the absolute value and the ^-component of the angular momentum have definite values (but the energy has no definite value). On the other hand, in the stationary states of an t Here and subsequently we ignore the (2/+l)-fold degeneracy with respect to the directions of the angular momentum relative to a fixed co-ordinate system. This degeneracy always occurs and is not physically important. If it is included, the total degree of degeneracy of the enenrv levels of a spherical top is (27+ 1) 2 . §103 Quantisation of the rotation of a rigid body 385 asymmetrical top the projection J^ of the angular momentum has, of course, no definite values, i.e. no definite values of k can be assigned to the energy levels. The operators ] ^ J v have matrix elements only for transitions in which k changes by unity, while J^ has only diagonal elements (see formulae (27.13), in which we must write /, k instead of L, M). Hence the operators J^ 2 , J 2 , /j 2 , and therefore fi, have matrix elements only for transitions with k -> k or &±2. The absence of matrix elements for transitions between states with even and odd k has the result that the secular equation of degree 2y + l immediately falls into two independent equations of degrees J and / + 1. One of these contains matrix elements for transitions between states with even k, and the other contains those for transitions between states with odd k. Each of these equations, in turn, can be reduced to two equations of lower degree. To do this, we must use the matrix elements defined, not With respect to the functions iff J]e , but with respect to the functions «A/fc + = (fo*+fo-*)/V 2 . / 103 7 n for = (fo*-fo-*)/\/2 (**0); the function ifi J0 is, of course, the same as ifi J0 +. Functions differing in the index + and — are of different symmetry (with respect to a reflection in a plane passing through the £-axis, which changes the sign of k), and hence the matrix elements for transitions between them vanish. Consequently we can form the secular equations separately for the + and — states. The Hamiltonian (103.1), and the commutation rules (103.3), have an unusual symmetry: they are invariant with respect to a simultaneous change in sign of any two of the operators J^J V , J v This symmetry formally corres- ponds to the group D 2 . Hence the levels of an asymmetrical top can be classified in accordance with the irreducible representations of this group. Thus there are four types of non-degenerate level, corresponding to the representations A, B ly B 2 , B z (see Table 7). It is easy to establish which states of the asymmetrical top belong to each of these types. To do so, we must find the symmetry properties of the func- tions (103.7), where the functions fo fc may be taken as simply the complex conjugate functions to the eigenfunctions of the angular momentum (which have the same symmetry; the complex conjugates are taken because of the change in sign on the right-hand sides of the commutation relations (103.3)) : where and <f> are spherical angles in £rj£-space. A rotation through an angle it about the £-axis (i.e. the symmetry operation C 2 (0 ) multiplies this function by (—1)*: C 2 <0: <A/*^(-l)*fo*. The operation C 2 ^> may be regarded as the result of successively per- forming an inversion and a reflection in the |£-plane; the first operation 386 Polyatomic Molecules §103 multiplies fok by (~l) J , and the second (a change in sign of <f>) is equivalent to changing the sign of k. Using the definition (28.6) of the functions ©j,-fc> we therefore have C 2 ( *>: <l>Jk->{-\) J+ Hj,-ic. Finally, the operation C 2 ® = C2<'>C 2 ^> gives Using these transformation rules, we find that the states corresponding to the functions (103.7) belong to the following types of symmetry: 4>Jk .+1 A B! B 2 \ $Jk < (103.8) ' J even, k even J even, k odd J odd, k even K J odd, k odd f J even, k even i?i / even, k odd 2?2 y odd, k even ^4 / odd, k odd U3 By simple counting it is easy to find the number of states of each type for a given value of/. The following numbers of states correspond to the types A and each of B h jB 2 , B3: A Bi, B2, B$ *7+i J even u (103.9) J odd ij-i |7+| Finally, we may add some remarks concerning the calculation of the matrix elements of various physical quantities characterising the top (or molecule). These are, of course, matrix elements with respect to the true rotational wave functions of the stationary states of the molecule (its electron state and vibrational state remaining unchanged). The wave functions of the symmetrical top are essentially the same as the angular part of the wave function of a diatomic molecule (with appropriate changes in the notation for the quantum numbers) : if the rotation is de- scribed by means of the Eulerian angles d, <f>, $ (Fig. 20, §58), the wave function of the state with quantum numbers k, J, Mj is fojMj = —e iM ->*eiW® lcJM (e), Ltt (103.10) where @ are the functions calculated in §82, Problem. The dependence of the matrix elements for a symmetrical top on the quantum numbers / and Mj is given by the formulae derived in §87 for a §103 Quantisation of the rotation of a rigid body 387 diatomic molecule (of zero spin). In formulae (87.1) — (87.4) K, Mr and A must be replaced respectively by the angular momentum / of the top, its projection Mj on the fixed .s-axis, and the projection k on the moving £-axis. The calculation of the matrix elements for the asymmetrical top is more complex. The solution of the secular equation derived from the matrix H J j\ gives linear combinations of the functions i/jjjc defined above which diagonalise the Hamiltonian. Replacing ifij^ in these combinations by the true wave functions (103.10) of the symmetrical top (with the same values of Mj in all these functions), we obtain the true wave functions of the stationary states of the asymmetrical top, described by values of the pair of numbers J, Mj. As a result, the calculation of the matrix elements of any quantity for the asymmetrical top reduces to that of the matrix elements for the sym- metrical top, which are already known. For the asymmetrical top there are selection rules for the matrix elements of transitions between states of the types A, B lt B 2 , B z \ these rules are easily obtained from symmetry considerations in the usual way. Thus, for the components of a vector physical quantity A we have the selection rules for ,4 g : A<->B 3 ®, B^^B^\ forA v : A++Bf», B^+^B^, for At : A<-> B^>, Bf*<-> B^. For clarity we show, as an index to the symbol for the representation, the axis about which a rotation has the character +1 in the representation concerned. PROBLEMS Problem 1. Calculate the matrix elements H J j\, for an asymmetrical top. Solution. From formulae (27.13) we find Ui%=Ur,% =*[/(/+ 1)"* 2 ]* : ( / 2)*+2 = 'ft+2 we 'Jc w ' 'Jfe+2 w)L - wr = -w>l = -ujr* for brevity, we everywhere omit the diagonal suffixes /, / of the matrix elements. Hence we have for the required matrix elements of H: 1 1 h*= ^(j+j^) U(J+i)-k*]+hWi2i c , H l + z = H T =i h2 (^-J^Vi(J-k)(J-k-l)(J+k+l)(J+k+2)]. 388 Polyatomic Molecules §103 The matrix elements with respect to the functions (103.7) are expressed in terms of the elements in formulae (1) by #£ = #1- =#*(*#!). &± = &+&,, &-=&-&, *+ 1+ (fc+2)+ (fc+2)- fc+2 V r h 2+ V 2 (2) PROBLEM 2. Determine the energy levels for an asymmetrical top with / = 1. Solution. The secular equation, of the third degree, falls into three linear equations. One of these is E x = H°+ whence E ± = W (i + r} (3) From this we can at once write down the other two energy levels, since it is obvious that the three moments of inertia I a, Ib, Ic enter the problem in a symmetrical manner. Hence it is sufficient simply to replace the moments of inertia I a, Ib once by 1 A , Ic and once by Ib, Ic- Thus E *= ih ih + r} E *= ih ir B + T} W The levels E u E 2 , E 3 belongf to the types B u B 2 , B 3 respectively, if I A , I B , I c are the moments of inertia about the $, 17, £ axes. Problem 3. The same as Problem 2, but for / = 2. Solution. The secular equation, of the fifth degree, falls into three linear equations and one quadratic. One of the linear equations is of the form E t = H2~, whence 2k 2 / 1 1 \ (5) a level of the type B x . Hence we at once conclude that there must be two other levels, of the types B 2 and B a : 2h 2 / 1 1 \ The equation of the second degree is 2h 2 / 1 1 \ E °=lu +ih iT B + T} H°+-E FP+ 0+ 0+ #2+ 0+ H*+-E 2+ = 0. (6) Solving this, we obtain ** - *(H + ^) ± ^[(^rt) , - 3 Gs; H 7i + i)]- (7) These levels belong to the type A. Problem 4. The same as Problem 2, but for / = 3. f This follows at once from considerations of symmetry. The energy E lt for instance, is symmetri- cal with respect to the moments of inertia I a and Ib, and this property belongs to the energy of a state whose symmetry about the £ and 17 axes is the same, i.e. a state of the type B\. §104 Interaction between the vibrations and the rotation of the molecule 389 Solution. The secular equation, of the seventh degree, falls into one linear equation and three quadratic. The linear equation is of the form E t = H%2, whence /J_ 1 1\ \T A + T B + TJ' E ^ 2h \T A + T B + T c y < 8) a level of the type A. One of the quadratic equations is equation (6) of Problem 3, with a different value of J. Its roots are 5h 2 /l 1\ h 2 ,r / 1 1\ 2 1 1 1 In E ™ ^ T\Ta + T b ) + T c ± h *J L\T A ~T B ) + ^IaTb'JaTc'JbTcS^ a level of the type B^. The remaining levels are obtained from these by permuting I a, Ib and Ic. Problem 5. Determine the splitting of the levels of a system having a quadrupole moment, in an arbitrary external electric field. Solution. Taking as co-ordinate axes the principal axes of the tensor d 2 <f>/8xi8xjc (see §75, Problem 4), we bring the quadrupole part of the Hamiltonian of the system to the form 8 = Ajj + Bjf + CjA A + B+C = 0. Owing to the complete formal analogy between this expression and the Hamiltonian (103.1), the problem under consideration is equivalent to that of finding the energy levels of an asymmetrical top, the only difference being that here the sum of the coefficients A + B + C = 0, and the angular momentum can have half-integral values also. For these the calculations must be done afresh by the same method, but for integral J we can use the results of Problems 2 to 4, obtaining the following values for the energy displacement AE for the first few values of J: 7=1: AE= -A,-B,-C; 7=3/2: AE = ± VP(^ 2 +5 2 +C 2 )/2]; 7=2: AE = 3A,3B,3C,±^/[6(A2 + B2 + C2)]. For J = 3/2 the energy levels remain doubly degenerate, in accordance with Kramers' theorem (§60). §104. The interaction between the vibrations and the rotation of the molecule Hitherto we have regarded the rotation and the vibrations as independent motions of the molecule. In reality, however, the simultaneous presence of both motions results in a peculiar interaction between them (E. Teller, L. Tisza and G. Placzek 1932-33). Let us start by considering linear polyatomic molecules. A linear molecule can execute vibrations of two types (see the end of §100): longitudinal vibra- tions with simple frequencies and transverse ones with double frequencies. We shall here be interested in the latter. A molecule executing transverse vibrations has in general some angular momentum. This is evident from simple mechanical considerations,-}- but it can also be shown by a quantum- mechanical discussion. The latter also enables us to determine the possible values of this angular momentum in a given vibrational state. f For example, two mutually perpendicular transverse vibrations with a phase difference of in can be regarded as a pure rotation of a bent molecule about a longitudinal axis. 390 Polyatomic Molecules §104 Let us suppose that some one double frequency a> a has been excited in the molecule. The energy level with the vibrational quantum number v a is (z; a +l)-fold degenerate. To this level there correspond the v a +l wave functions Kdfa, = constant x e-c^Q^+Q.^H.J^Q^H^^Q^), where ^ a i+^ a2 = ^ a > or any independent linear combinations of them. The total degree (in Q al and Q a2 together) of the polynomial by which the exponential factor is multiplied is the same in all these functions, and is equal to v a . It is evident that we can always take, as the fundamental functions, linear combinations of the functions ift v v of the form ^ai^ai *,*. = constant x e -ca'(Q ai J +Q a , 8 )/2 [{Q al +iQ^r^)l^Q^-iQ^^)l2 + ... j. (104.1) The square brackets contain a determinate polynomial, of which we have written out only the highest term. / a is an integer, which can take the v a -\-l different values v a , v a —2,v (X —4;..., —v^. The normal co-ordinates Q al , Q a2 of the transverse vibration are two mutually perpendicular displacements off the axis of the molecule. Under a rotation through an angle cf> about this axis, the highest term of the poly- nomial (and therefore the whole function ip v z ) is multiplied by gi0(i> a +7 a )/2£-i<Mv a -Z a )/2 _ gi* a #. Hence we see that the function (104.1) corresponds to a state with angular momentum / a about the axis. Thus we reach the result that, in a state where the double frequency a> a is excited (with quantum number v a ), the molecule has an angular momentum (about its axis) which takes the values l« = »«, ^~ 2 > ®«-4» - . -»«• (104.2) This is called the vibrational angular momentum of the molecule. If several transverse vibrations are excited simultaneously, the total vibrational angular momentum is equal to the sum S/ a . On being added to the electron orbital angular momentum, it gives the total angular momentum / of the molecule about its axis. The total angular momentum / of the molecule cannot be less than the angular momentum about the axis (just as in a diatomic molecule), i.e. J takes the values In other words, there are no states with J = 0, 1, ... , |/| — 1. For harmonic vibrations, the energy depends only on the numbers v^ and not on l a . The degeneracy of the vibrational levels (with respect to the §104 Interaction between the vibrations and the rotation of the molecule 391 values of / a ) is removed by the presence of anharmonic vibrations. The removal is not complete, however: the levels remain doubly degenerate, the same energy belonging to states differing by a simultaneous change of sign of all the / a and of /. In the next approximation (after that of harmonic motion), a term quadratic in the angular momenta l a , of the form S g a /J fi (the g ufi being constants), appears in the energy. This remaining double degeneracy is removed by an effect similar to the A-doubling in diatomic molecules. When we turn to non-linear molecules, we must first of all make the following remark, which has a purely mechanical significance. For an arbi- trary (non-linear) system of particles, the question arises how we can at all separate the vibrational motion from the rotation; in other words, what we are to understand by a "non-rotating system". At first sight it might be thought that the vanishing of the angular momentum, S mrxv = (104.3) (the summation being over the particles in the system), could serve as a criterion of the absence of rotation. However, the expression on the left- hand side is not the complete derivative, with respect to time, of any function of the co-ordinates. Hence the above equation cannot be integrated with respect to time in such a way as to be formulated as the vanishing of some function of the co-ordinates. This, however, is necessary if a reasonable definition of the concepts of "pure vibrations" and "pure rotation" is to be possible. As a definition of the absence of rotation, we must therefore use the condition S mr xv = 0, (104.4) where r are the radius vectors of the equilibrium positions of the particles. Putting r = r +u, where u are the displacements in small vibrations, we have v = f = u. The equation (104.4) can be integrated with respect to time, giving 2wr xu=0. (104.5) The motion of the molecule will be regarded as a combination of the purely vibrational motion, in which the condition (104.5) is satisfied, and the rotation of the molecule as a whole.f Writing the angular momentum in the form 2 mrxv = S mr Xv+ 2 muxv, we see that, in accordance with the definition (104.4) of the absence of rotation, the vibrational angular momentum must be understood as the sum Smuxv. However, it must be borne in mind that this angular momentum, f The translational motion is supposed removed from the start, by choosing a system of co-ordinates in which the centre of mass of the molecule is at rest. 392 Polyatomic Molecules §104 being only a part of the total angular momentum of the system, is not con- served. Hence only a mean value of the vibrational angular momentum can be ascribed to each vibrational state. Molecules having no axis of symmetry of order above the second belong to the asymmetrical-top type. In a molecule of this type, all the frequencies are simple (their symmetry groups have only one-dimensional irreducible representations). Hence none of the vibrational levels is degenerate. In any non-degenerate state, however, the mean angular momentum vanishes (see §26). Thus, in a molecule of the asymmetrical-top type, the mean vibra- tional angular momentum vanishes in every state. If, among the symmetry elements of the molecule, there is one axis of order higher than the second, the molecule is of the symmetrical-top type. Such a molecule has vibrations with both simple and double frequencies. The mean vibrational angular momentum of the former again vanishes. To the double frequencies, however, there corresponds a non-zero mean angular momentum component along the axis of the molecule. It is easy to find an expression for the energy of the rotational motion of the molecule (of the symmetrical-top type), taking into account the rotational angular momentum. The operator of this energy differs from (103.5) in that the rotational angular momentum of the top is replaced by the difference between the total (conserved) angular momentum J of the molecule and its vibrational angular momentum J(») : tfrot = — — (J- K> 2 +P 2 (^ i-XJs- Ja 2 - (104.6) The required energy is the mean value H TO t. The terms in (104.6) which contain the squared components of J give a purely rotational energy which is the same as (103.6). The terms which contain the squared components of J(f) give constants independent of the rotational quantum numbers and may be omitted ; the terms which contain products of components of J and JW constitute the interaction here considered between the vibrations of the molecule and its rotation. This is called the Coriolis interaction (since it corresponds to the Coriolis forces in classical mechanics). In averaging these terms it must be borne in mind that the mean values of the transverse (£,17) components of the vibrational angular momentum are zero. The energy of the Coriolis interaction is therefore Ecox=-h*kk v IIc, (104.7) where the integer k is, as in §103, the component of the total angular momen- tum along the axis of the molecule, and k v = 1^ is the mean value of the component of the vibrational angular momentum for the vibrational state concerned; k v , unlike k, is not an integer. Finally, let us consider molecules of the spherical-top type. These include molecules whose symmetry is that of any of the cubic groups. Such molecules §104 Interaction between the vibrations and the rotation of the molecule 393 have simple, double and triple frequencies (there being one-, two- and three- dimensional irreducible representations of the cubic groups). The degeneracy of the vibrational levels is, as usual, partly removed by the presence of anharmonic motion; when these effects have been taken into account there remain, apart from the non-degenerate levels, only doubly and triply degene- rate levels. Here we shall discuss these levels that are split by the presence of anharmonic motion. It is easy to see that, for molecules of the spherical-top type, the mean vibrational angular momentum is zero not only in the non- degenerate vibrational states but also in the doubly degenerate ones. This follows from simple considerations based on symmetry properties. The mean angular momentum vectors in two states belonging to the same degenerate energy level must be transformed into each other in all possible symmetry trans- formations of the molecule. None of the cubic symmetry groups, however, allows the existence of two directions transformed only into each other; only sets of three or more directions are so transformed. From these arguments it follows that, in states corresponding to triply degenerate vibrational levels, the mean vibrational angular momentum is non-zero. After averaging over the vibrational state, this angular momentum is represented by an operator whose matrix elements correspond to transitions between three mutually degenerate states. In accordance with the number of these states, this operator must have the form £l, where 1 is the operator of an angular momentum of unity (for which 2/+ 1 = 3) and £ is a constant characterising the vibrational level in question. The Hamiltonian of the rotational motion of the molecule is tfrot^WXJ-H and, after averaging, becomes the operator fc2 fe2-^ — ffi # rot = -J2+ J<*>2_ £2j.i. (104.8) 21 21 I The eigenvalues of the first term give the ordinary rotational energy (103.4); the second term gives an unimportant constant, which does not depend on the rotational quantum number. The last term in (104.8) gives the desired energy of the Coriolis splitting of the vibrational level. The eigenvalues of the quantity J . 1 are calculated in the usual way ; it has (for a given J) three different values corresponding to the values /+1, /— 1, / of the vector J + l. The result is EcoT iJ+1) =-hHJII, ^ j? Cor w-i) = hH(J+ 1)//, (104.9) Ecot (J) =h%II. J 394 Polyatomic Molecules §105 §105. The classification of molecular terms The wave function of a molecule is the product of the electron wave func- tion, the wave function of the vibrational motion of the nuclei, and the rota- tional wave function. We have already discussed the classification and types of symmetry of these functions separately. It now remains for us to examine the question of the classification of molecular terms as a whole, i.e. of the possible symmetry of the total wave function. It is clear that, if the symmetry of all three factors with respect to some transformation is given, the symmetry of the product with respect to that transformation is determined. For a complete description of the symmetry of the state, we must also specify the behaviour of the total wave function when the co-ordinates of all the particles in the molecule (electrons and nuclei) are inverted simultaneously. The state is said to be negative or positive, according as the wave function does or does not change sign under this transformation.-}- It must be remembered, however, that the characterisation of the state with respect to inversion is significant only for molecules which do not possess stereoisomers. If stereoisomerism is present, the molecule assumes on inversion a configuration which can by no rotation in space be made to coincide with the original configuration; these are the "right-hand" and "left-hand" modifications of the substance.J Hence, when stereoisomerism is present, the wave functions obtained from each other on inversion belong essentially to different molecules, and it is meaningless to compare them. || We have seen in §86 that, for diatomic molecules, the spin of the nuclei exerts an important indirect effect on the arrangement of the molecular terms by determining their degree of degeneracy, and in some cases entirely forbidding levels of a certain symmetry. The same is true for polyatomic molecules. Here, however, the investigation of the problem is considerably more complex, and requires the application of the methods of group theory to each particular case. The idea of the method is as follows. The total wave function must con- tain, besides the co-ordinate part (the only part we have considered so far), a spin factor, which is a function of the projections of the spins of all the nuclei on some chosen direction in space. The projection a of the spin of a nucleus takes 2/+1 values (where i is the spin of the nucleus); by giving to all the oj, a 2 , ... , a N (where N is the number of atoms in the molecule) all possible values, we obtain altogether (2^ + 1) (2/ 2 +l) ... (2^ + 1) different values of the spin factor. In each symmetry transformation, certain nuclei (of the same kind) change places, and if we imagine the spin values to "remain fixed", the transformation is equivalent to an interchange of spin values among t We use the same customary, though unfortunate, terminology as for diatomic molecules (§86). % For stereoisomerism to be possible, the molecule must have no symmetry element pertaining to reflection (i.e. no centre of symmetry, plane of symmetry, or rotary-reflection axis). || Strictly speaking, quantum mechanics always gives a non-zero probability for the transition from one modification to the other. This probability, however, which relates to the passage of nuclei through a barrier, is so small that the phenomenon can always be neglected. §105 The classification of molecular terms 395 the nuclei. Accordingly, the various spin factors will be transformed into linear combinations of one another, thus giving some representation (in general reducible) of the symmetry group of the molecule. Decomposing this into irreducible parts, we find the possible types of symmetry for the spin wave function. A general formula can easily be written down for the characters x sp (^) of the representation given by the spin factors. To do this, it is sufficient to notice that, in a transformation, only those spin factors are unchanged in which the nuclei changing places have the same a a \ otherwise, one spin factor changes into another and contributes nothing to the character. Bearing in mind that a a takes 24+1 values, we find that Xs P (G0=II(2* a +l), (105.1) where the product is taken over the groups of atoms which change places under the transformation G considered (there being one factor in the product from each group). We are, however, interested not so much in the symmetry of the spin function as in that of the co-ordinate wave function (by which we mean its symmetry with respect to interchanges of the co-ordinates of the nuclei, the co-ordinates of the electrons remaining unchanged). These two symmetries are directly related, however, since the total wave function must remain unchanged or change sign when any pair of nuclei are interchanged, accord- ing as they obey Bose statistics or Fermi statistics (in other words, it must be multiplied by ( — 1) 2 \ where i is the spin of the nuclei that are interchanged). Introducing the appropriate factor in the characters (105.1), we obtain the system of characters x(G) for the representation containing all the irreducible representations by which the co-ordinate wave functions are transformed: X (G) = n(2f.+ l)(-l)^».-« (105.2) where n a is the number of nuclei in each group which change places under the transformation in question. Decomposing this representation into irreducible parts, We obtain the possible types of symmetry of the co-ordinate wave functions of the molecule, together with the degrees of degeneracy of the corresponding energy levels (here and later we mean the degeneracy with respect to the different spin states of the system of nuclei).f Each type of symmetry of the states is related to definite values of the total spins of the groups of equivalent nuclei in the molecule (i.e. groups of nuclei which change places under the various symmetry transformations of the molecule). This relation is not one-to-one; each type of symmetry of states can, in general, be brought about with various values of the spins of equivalent groups. The relation can also be established, in any particular case, by means of group theory. t The degree of degeneracy of the level in this respect is often called its nuclear statistical weight; see the last footnote to §86. 396 Polyatomic Molecules §105 As an example, let us consider a molecule of the asymmetrical-top type, the ethylene molecule C^H^ (Fig. 43g), with the symmetry group D 2h . The index to the chemical symbol indicates the isotope to which the nucleus belongs; this indication is necessary, since the nuclei of different isotopes have different spins. In this case, the spin of the H 1 nucleus is \, while the C 12 nucleus has no spin. Hence we need consider only the hydrogen atoms. We take the system of co-ordinates shown in Fig. 43g; the z-zxis is per- pendicular to the plane of the molecule, while the #-axis is along the axis of the molecule. A reflection in the ry-plane leaves all the atoms fixed, while other reflections and rotations interchange the hydrogen atoms in pairs. From formula (105.2) we have the following characters of the represen- tation : E <j(xy) o(xz) a{yz) I C 2 (x) C 2 (y) C 2 (z) 16 16 4 4 4 4 4 4 Decomposing this representation into irreducible parts, we find that it con- tains the following irreducible representations of the group D ih : 7A g , 3B lg , 3B 2u , 3B 3u . The figures show the number of times each irreducible represen- tation appears in the reducible one; these numbers are also the nuclear statistical weights of the levels with the corresponding symmetry.f The classification of the states of the ethylene molecule thus obtained relates to the symmetry of the total (co-ordinate) wave function, including the electron, vibrational and rotational parts. Usually, however, it is of interest to arrive at these results from a different point of view. Knowing the possible symmetries of the total wave function, we can find at once which rotational levels are possible (and with what statistical weights) for any prescribed electron and vibrational state. Let us consider, for instance, the rotational structure of the lowest vibra- tional level (that for which the vibrations are not excited at all) of the normal electron term, assuming that the electron wave function of the normal state is completely symmetrical (as is the case for practically all polyatomic molecules). Then the symmetry of the total wave function with respect to rotations about the axes of symmetry is the same as the symmetry of the rotational wave function. Comparing this with the results obtained above, we therefore conclude that in the ethylene molecule the rotational levels of the types A and B ± (see §103) are positive with statistical weights 7 and 3, while those of the types B 2 and B3 are negative with statistical weight 3. As with diatomic molecules (see the end of §86), owing to the extreme weakness of the interaction between the nuclear spins and the electrons, transitions between states of different nuclear symmetry in the ethylene molecule do not occur in practice. Hence molecules in such states behave like different modifications of the substance. Thus ethylene C 12 2 H 1 4 has four modifications, with nuclear statistical weights 7, 3, 3, 3. t The relation between the symmetry of states and the values of the total spin of the four hydrogen nuclei in the ethylene molecule is derived in Problem 1. §105 The classification of molecular terms 397 In reaching this conclusion it is important that states with different symmetry belong to different energy levels (the intervals between which are large compared with the interaction energy of nuclear spins). The conclusion is therefore invalid for molecules in which there exist states of different nuclear symmetry belonging to the same degenerate energy level. Let us consider another example, the ammonia molecule N 14 H 1 3 , of the symmetrical-top type (Fig. 41), whose symmetry group is C 3v . The spin of the nucleus N 14 is 1, and that of H 1 is £. Using formula (105.2), we find the characters of the representation of the group C 3v in which we are interested: E 2Cq 3 <7„ 24 6 -12 It contains the following irreducible representations of the group C 3v : 12A 2 , 6E. Thus two types of level are possible; their nuclear statistical weights aref 12 and 6. The rotational levels of a symmetrical top are classified (for a given /) according to the values of the quantum number k. Let us consider, as in the previous example, the rotational structure of the normal electron and vibrational state of the NH 3 molecule (i.e. we suppose the electron and vibra- tional wave functions to be completely symmetrical). In determining the symmetry of the rotational wave function, we must bear in mind that it is meaningful to speak of its behaviour only with respect to rotations about axes. Hence we replace the planes of symmetry by axes of symmetry of the second order perpendicular to them, a reflection in a plane being equivalent to a rotation about such an axis, followed by an inversion. In the present case, therefore, instead of the group C 3v we have to consider the isomorphous point group D 3 . The rotational wave functions with k = ±\k\ are multiplied by e ±2lTiW / z respectively under a rotation C 3 about a vertical axis of the third order, while under a rotation U 2 about a horizontal axis of the second order they change into each other, thus giving a two-dimensional representation of the group D s . If \k\ is not a multiple of three, this representation is irreduc- ible; it is E. The representation of the group C 3V corresponding to the total wave function is obtained by multiplying the character x(U 2 ) by 1 or — 1, according as the term is positive or negative. Since, however, in the repre- sentation E we have x(^) = 0> we obtain the same representation E in either case (but this time as a representation of the group C 3v , and not D 3 ). Bearing in mind the results obtained above, we thus conclude that, when \k\ is not a multiple of three, both positive and negative levels are possible, with nuclear statistical weights of 6 (the symmetry of the total co-ordinate wave function being of the type E). When \k\ is a multiple of three (but not zero), the rotational functions t A total spin of the hydrogen nuclei of 3/2 corresponds to the terms of symmetry A 2 , and one of 1/2 to those of symmetry E. 398 Polyatomic Molecules §105 give a representation (of the group D 3 ) with characters E 2C 3 3U 2 2 2 This representation is reducible, and divides into the representations A x , A%. In order that the total wave function should belong to the representation A 2 of the group C 3 „, the rotational level A x must be negative and A % positive. Thus, when \k\ is a multiple of three and not zero, both positive and negative levels are possible, with nuclear statistical weights of 12 (levels of the type A 2 ). Finally, only one rotational function corresponds to an angular momentum component k = 0; it gives a representation with charactersf E 2C, 3U 9 1 1 (~iy If the total wave function has the symmetry A 2 , its behaviour with respect to inversion must therefore be given by the factor ( — 1) J+1 . Thus, for k = 0, levels with even and odd / can only be negative and positive respectively; the statistical weight is 6 in either case (levels of the type A 2 ). Summarising these results, we have the following table of possible states for various values of the quantum number k for the normal electron and vibrational term of the molecule N 14 H 1 3 (the symbols + and - denoting positive and negative states). + - | k | not a multiple of 3 6E 6E | A | a multiple of 3 12^4 2 12^ 2 h _ n / J even _ 6A * * ~ U \ J odd 6A 2 For given J and k, the energy levels of the NH3 molecule are in general degenerate (see also the table for ND 3 in Problem 3). This degeneracy is partly removed by a peculiar effect due to the flat shape of the ammonia molecule and the small mass of the hydrogen atoms. By a fairly small vertical displacement of the atoms in this molecule a transition can be brought about between two configurations obtained from one another by a reflection in a plane parallel to the base of the pyramid (Fig. 44). These transitions cause a splitting of the levels, separating positive and negative levels (an effect similar to the one- dimensional case considered in §50, Problem 3). The magnitude of the splitting is proportional to the probability of passage of the atoms through the "potential barrier" separating the two configurations of the molecule. Although this probability is comparatively high in the ammonia t On rotation through an angle w, the eigenfunction of the angular momentum with magnitude J and component zero is multiplied by (— \) J . §105 The classification of molecular terms 399 molecule, owing to the above-mentioned properties, the splitting is still small (1 x 10- 4 eV). An example of a molecule of the spherical-top type is discussed in Problem 5. C2(y) C 2 (x) 1 1 2 2 PROBLEMS Problem 1. Find the relation between the symmetry of the state of the C 12 2H 1 4 molecule and the total spin of the hydrogen nuclei in the molecule. Solution^ The total spin of the four H 1 nuclei can take the values / = 2, 1, 0, and its component Mi takes values from 2 to —2. Let us consider the representations given by the spin factors for each value of Mi, beginning with the largest. The value Mi = 2 corresponds to only one spin factor, in which all the nuclei have a spin component + i- The value Mi = 1 corresponds to four different spin factors differing as regards the nucleus which has spin component -£. Finally, the value Mi = is given by six spin factors, depending on the pair of nuclei which have spin components — \. The characters of the three corresponding representations are as follows: E a(xy) a(xz) o(yz) I Cz(z) Mi = 2 1 1 1 1 1 1 Mi = \ 4 4 Mi = 6 6 2 2 2 2 The first of these representations is the unit representation A g ; since the value Mi = 2 can occur only for J = 2, we conclude that a state with symmetry A g corresponds to spin 1 = 2. The value Mi = 1 can occur for both / = 1 and / = 2. Subtracting the first representa- tion from the second and decomposing the result into irreducible parts, we find that states Big, Bz u , Bzu correspond to spin 1 = 1. Finally, the value Mi = can occur in all cases where Mi = 1 is possible, and also for 1 = 0. Subtracting the second representation from the third, we find two states Ag corres- ponding to spin 1 = 0. Problem 2. Determine the types of symmetry of the total (co-ordinate) wave functions, and the statistical weights of the corresponding levels, for the molecules C 12 2 H%, C'^H 1 ^ N 14 2 14 4 (all these molecules are of the same form; the spins are i(H*) = 1, i(C 13 ) = £, t(N 14 ) = 1). Solution. By the method shown in the text for the molecule C^H^, we find the follow- ing states (the axes of co-ordinates being taken the same as above) : Molecule + — C 12 2 H 2 4 27 A g , lSBig lSBzu, WBau Ci3 2 H 1 4 16A g , \2Big \2Bzu, 2\Bzu Ni4 2 oi 6 4 6A g 3i?3u f A method of solving problems of this kind, based on the theory of permutation groups, is given by E. G. Kaplan, Soviet Physics JETP 37(10), 747, 1960. 400 Polyatomic Molecules §105 Problem 3. The same as Problem 2, but for the molecule N 14 H a 3 . Solution. In the way shown in the text for the molecule N 14 !! 1 ., we find the states 30.4,, 3A 2 , 2AE. In the normal electron and vibrational state, the following terms are possible for various values of the quantum number k : \k\ not a multiple of 3 \k\ a multiple of 3 * = «{ JST Problem 4. The same as Problem 2, but for the molecule C^H 1 , (see Fig. 43f; the symmetry is D a d ). Solution. The possible states are of the types 7A ig , 1A 1U , 3A ig , 13A iV , 9E g , 11E U . In the normal electron and vibrational state, the following levels are obtained: -r- 24£ 2AE 30A U 3A 2 30A lt 3A % 30A t 3A 2 3A % 30A t + — \k\ not a multiple of 3 9E g UE U \k\ a multiple of 3 * A\ g , 3A ig 1A 1U , 13A, 4 j J even i /odd 7A ig \A\u 3A 2g 13 A%u k = Problem 5. The same as Problem 2, but for the methane molecule C^Wi (the C atom is at the centre of a tetrahedron with the H atoms at the vertices). Solution. The molecule is of the spherical-top type, and has the symmetry T&- Follow- ing the same method, we find that the possible states are of the types 5Az, IE, 3Fi (cor- responding to a total spin of the molecule of 2, 0, 1 respectively). The rotational states of a spherical top are classified according to the values of the total angular momentum /. The 2/4-1 rotational functions belonging to a particular value of / give a (2/ + l)-dimensional representation of the group O, which is isomorphous with the group Tdl it is obtained from the latter by replacing all planes of symmetry by axes of the second order perpendicular to them. The characters in this representation are given by formula (98.1). Thus for example, for J = 3 we obtain a representation with characters E 8C 3 6C 2 6C 4 3C 4 2 1 -1 _i _i This contains the following irreducible representations of the group O: A if F u F 2 . Again considering the rotational structure of the normal electron and vibrational term, we therefore conclude that, for J = 3, the states with a symmetry Az of the total wave function can only be positive, while those of type F x can be either positive or negative. For the first few values of/ we thus obtain the following states (which we write together with their statistical weights) : + - / = — 5A 2 /=1 / = 2 / = 3 / = 4 IE IE, 3F X 5A 2 , 3F t 3F X IE, 3F X SA 2 , IE, 3F X CHAPTER XIV ADDITION OF ANGULAR MOMENTA §106. 3/-symbols The rule of addition of angular momenta deduced in §31 gives the possible values of the total angular momenta of a system consisting of two particles (or more complex components) with angular momenta j\ and j%.-\ This rule is in fact closely related to the properties of wave functions with respect to spatial rotations, and follows immediately from the properties of spinors considered in Chapter VIII. The wave functions of particles with angular momenta j± and j% are sym- metrical spinors of ranks 2/i and 2/2, and the wave function of the system is their product, 0(i) ht^mH^-. (106.1) 2/1 2/2 Symmetrising this product with respect to all the indices, we obtain a sym- metrical spinor of rank 2(/i +72), corresponding to a state with total angular momentum 71 +7*2. If we contract the product (106.1) with respect to one pair of indices, of which one must belong to if/W and the other to tpW (since otherwise the result is zero), the symmetry of each of the spinors j/t* 1 ) and «/r( 2 > shows that it does not matter which indices are taken from A, /*, ... and p, cr, ... . After symmetrisation we obtain a symmetrical spinor of rank 2(/i+/2—l), corresponding to a state with angular momentum j\ +J2— 1 •% Continuing this process, we find, in agreement with the rule already known, that/ takes values from j\ +j% to |/i— /2I, each occurring once. For a complete solution of the problem of the addition of angular momenta, we must also consider the problem of constructing the wave function of a system with a given total angular momentum from those of its two component particles. t Strictly speaking, we shall always be considering (without explicitly mentioning the fact each time) a system whose parts interact so weakly that their angular momenta may be regarded as con- served in a first approximation. All the results given below apply, of course, not only to the addition of the total angular momenta of two particles (or systems) but also to the addition of the orbital angular momentum and spin of the same system, assuming that the spin-orbit coupling is sufficiently weak. % To avoid misunderstanding, the following comment is useful. The wave function of a system of two particles is always a spinor of rank 2{ji-\-J2), and this is in general not equal to 2j, where / is the total angular momentum of the system. Such a spinor may, however, be equivalent to a spinor of lower rank. For example, the wave function of a system of two particles with angular momenta / 1 =ji = £ is a spinor of rank two ; but if the total angular momentum j = 0, this spinor is anti- symmetrical, and therefore reduces to a scalar. In general, the total angular momentum / determines the symmetry of the spinor wave function of the system : this is symmetrical with respect to 2j indices and antisymmetrical with respect to the remainder. 401 402 Addition of Angular Momenta §106 Let us begin with the simple case of the addition of two angular momenta to give a zero total angular momentum. Here we must evidently have /i = j2 and angular momentum components m\ = — m^. Let tftj m be the normalised wave functions of the states of one particle with angular momentum j and component thereof m (in the non-spinor representation). The required wave function To of the system is the sum of the products of the wave functions of the two particles with opposite values of m : T = 2 (-iy- m (1 W (2) y,-m, (106.2) where/ is the common value of/i and 72- The factor preceding the sum is due to the normalisation. The coefficients in the sum must all have the same absolute value, since all values of the components m of the angular momenta of the particles must be equally probable. The sequence of signs in (106.2) is easily found by means of the spinor representation of the wave functions. In spinor notation the sum in (106.2) is a scalar (the total angular momentum of the system being zero) ^•"^V..* (106.3) formed from two spinors of rank 2/. Using this, we find the signs in (106.2) directly from (57.3). It should be borne in mind, however, that in general only the relative signs of the terms in the sum (106.2) are determinate, while the sign of the whole sum may depend on the "order of addition" of the angular momenta. For, if we lower all spinor indices (j+m ones and/— m twos) in j/^ 1 ) and raise them in j/»( 2 >, the scalar (106.3) is multiplied by (— l) 2 ^, and therefore changes sign when/ is half-integral. Next we consider a system with zero total angular momentum consisting of three particles with angular momenta /1, j'2, 7*3 and components thereof mi, m2, m%. The condition for the total angular momentum to be zero is that m>\ + W2 + m% = and j\, j'2, js have values such that each of them can be obtained by vector addition of the other two, i.e. geometrically /1, j'2, jz must be the sides of a closed triangle. In other words, each of them lies between the difference and the sum of the other two : |/i -/2I </3 <ji+J2, etc. It is evident that the algebraic sum /1 4-/2 +/s is an integer. The wave function of the system under consideration is the sum Wo= 2 ( Jl J2 ^VW^m^W (106.4) •*— ' \7Wi 7W2 mzt WJl,Wl2,Wl3 taken over the values of each rm from — j\ toji. The coefficients in this formula §106 3j-symbols 403 are termed Wigner 3j-symbols. By definition they are non-zero only if W1 + W2 + W3 = 0. When the suffixes 1, 2, 3 are permuted, the wave function (106.4) can change only by an unimportant phase factor. The 3j-symbols can in fact be defined as purely real quantities (see below), and then the indeterminacy of To can consist only in its sign as a whole being indefinite (as is true of the function (106.2) also). This means that interchanging the columns of a 3/-symbol can either leave it unchanged or change its sign. The most symmetrical way of defining the coefficients in the sum (106.4), which is the definition generally used for the 3/-symbols, is as follows. In spinor notation, T is a scalar formed by contracting the product of the three spinors j/fW^-, 0<2)*/«- f 0(3) V... w ith respect to all pairs of indices belonging to two different spinors. In each pair belonging to particles 1 and 2 the spinor index will be written superior with ifjM and inferior with 0< 2 ) ; in a pair belong- ing to particles 2 and 3, superior with j/»< 2 > and inferior with </f< 3 > ; and in a pair belonging to particles 3 and 1, superior with ifj® and inferior with 0W. It is easily seen that the total number of pairs of each kind is/i +72 —7*3,72 +73 —ji, 71+73—72 respectively. This rule determines uniquely the sign of To. It is evident that, with this definition, cyclic interchange of the indices 1, 2 and 3 leaves T unchanged. This means that the 37-symbol is unchanged when its columns are cyclically permuted. Interchange of any two indices is easily seen to require the raising of the lower indices and lowering of the upper indices in all 71 +72 +73 pairs. This means that T is multiplied by (_l)A+/2+i 3 ; in other words, the 37-symbols have the property lh h h\ _ ( _ l)h+h+l ,(h h h\ etc _ (1065) W2 mi mz/ wi mz m 3 ) i.e. they change sign when two columns are interchanged if7i+72+73 is odd. Finally, we easily see that / h h h\ { _ 1)h+M /h h h\ {106 . 6) \ — mi —mi —mz) Wi m% mz) a change in the sign of the z- component of each angular momentum can be regarded as the result of a rotation through an angle -n- about the jy-axis, and this is equivalent to raising all the lower spinor indices and lowering all the upper ones (see (58.11)). From (106.4) we can derive an important formula which gives the wave function Yj m of a system consisting of two particles and having given values of 7 and m. To do so, we consider the particles 1 and 2 together as one system. Since the angular momentum j of this system together with the angular momentum J3 of particle 3 gives a total angular momentum of zero, we must have/ = 73, m = —mz. According to (106.2) we can then write To = J- ix 2 ( - 1 )'- w, *W (8) /. -m- (106.7) 14 404 Addition of Angular Momenta §106 This formula is to be compared with (106.4) (in which we replace jz, mz by /, — m). Here, however, we must first take into account the fact that the rule for constructing the sum in (106.7) according to (106.3) does not corres- pond to the rule for constructing the sum (106.4): to bring (106.7) to the form (106.4) we must, as is easily seen, interchange pairs of upper and lower indices corresponding to particles 1 and 3. This leads to an additional factor (-\)h-h+h. The result is W jm = (-1)W.+«V(2/ + 1) y ( Jl 32 J VW"W (106.8) *— ' wi t»2 — tn/ where the summation over m\ and m^ is subject to the condition m± + mi = m. Formula (106.8) gives the required expression for obtaining the wave function of a system from those of its two particles, which have definite angular momenta j\ and }<i> It can be written in the form Vjm= 2 0m tt) W (2) 'a»»2 (m 2 = m-m 1 ). (106.9) mum* numa The coefficients! c L m w =(-iy^^v(2y+i)(- /1 n J ) (106.10) form the matrix of the transformation from the complete orthonormal set of (2/i+l)(2/2+l) wave functions of states with definite mi, mi to the similar set with definite/, m (for given values oijijjz). As we know (see §12), such a matrix is unitary. Hence we can immediately write down the inverse trans- formation: h+it ^W^y.-, = 2 cs 'T mzX¥ ^^^> ( 106 - n ) i=\h-h\ mimz where we have also used the fact that the coefficients C are real. According to the general rules of quantum mechanics, the squares of the coefficients in the expansion (106.11) give the probability for the system to have any par- ticular value of/ (for given /i, mi and 72, mi). The unitarity of the transformation (106.9) means that its coefficients satisfy certain orthogonality conditions. According to formulae (12.5) and (12.6) 2 0m 0'm' miwi2 minis mi,t»2 t Called vector addition coefficients or Clebsch-Gordan coefficients. The notation Cj m,) 2 m 2 or (jlJ2J T n\jljzmim2) is also used in the literature. §106 3j-symbols 405 - (2;+i) 2 ( h h j )( hh f ) *—> \mi m 2 — ml \m\ m 2 —m I 2CJm Qim miffls mi'rm j,m 2mv( hh i )( h . h . *—< \m± m 2 —ml \m± m% — ml = 8 8 . (106.13) JBlWll' 77127712' The explicit general form of the 3j- symbols is quite lengthy. It can be written asf (h h h\ _ r O'i +h -h) l -(ji ~h +js)i( -ji +h +73)1 -1 1/2 Vmi m2 m3/ L ( J i -h/2 -+-J3 + 1 ) - J x [(ji + wi)!(ji - mi)!(j 2 + m 2 )!(j2 - m 2 )\(js + m s)Kh ~ »*3)!] 1/2 x ( — \\z+h-h-m* *!(ji +72 -;s - *)!(ji - w i - *)Ki2 + m 2 - ar)!(j3 -72 + «i + *)!(js -/1 - ^ 2 +*) ! (106.14) The summation is over all integers z but, since the factorial of a negative number is infinite, the sum contains only a finite number of terms. The coefficient of the sum is obviously symmetrical in the suffixes 1, 2, 3; the symmetry of the sum itself appears if the values of the summation variable z are interchanged. Besides the symmetry properties (106.5) and (106.6), which follow imme- diately from the definition of the 3/-symbols, the latter also have other symmetry properties, though the derivation of these is more complex and will not be given here.J The properties in question can be conveniently formulated in terms of a three-by-three array of numbers derived from the parameters of the 3/-symbol as follows : . . p/2 +73-.71 73+71-72 71 + 72-73-1 ( ) = I h— m i j'2-m h—mz I; (106.15 \mi m 2 mz I I I I— j'i 4- m-t Jo 4- mo 11 4- m* -■* 7l + »*l 72 + ^2 73 + W3 t The coefficients in (106.9) were first calculated by E. Wigner (1931). Their symmetry properties and the symmetrical expression (106.14) were first derived by G. Racah (1942). The most direct method of calculation is probably to go immediately from the spinor representation of *Fo (approp- riately normalised) to the representation in the form of the sum (106.4) by means of the correspon- dence formula (58.1); it may be noted that, since the coefficient in this formula is real, so also must be the 3/-symbols. Another derivation is given by A. R. Edmonds, Angular Momentum in Quantum Mechanics, Princeton 1957. The table of 3/-symbols given below is also taken from Edmonds' book. { See T. Regge, 72 nuovo cimento [10] 10, 544, 1958. 406 Addition of Angular Momenta §106 the sum of the numbers in each row and each column of this array is/i +72 +73« Then (1) interchange of any two columns of the array multiplies the 3/- symbol by (—\)h+U+h (the same property as that given by (106.5)); (2) the same is true for interchange of any two rows (for the two lower rows, the same property as that given by (106.6)); (3) the 3/-symbol is unchanged when the rows and columns of the array are interchanged. Some of the simpler formulae for particular cases will be given here. The value (3 3 °\ 1 r i = (-iy-» (106.16) corresponds to formula (106.2). The formulae fh h h+j h/2 \ j = (-i)h-h [(2ji)\(2J2)l(ii+J2 + mi + m 2 )Kh+3'2-mi-m 2 )\ ~| 1/2 ^ .„. , (106.17) (2/i + 2j 2 + l)!(ji + wn)!(ji - mi)!(; 2 + m 2 )!0' 2 - ^2)U V/i — 71 — ms mzl [(2/i) ! ( -h +32 +J3)Kh +h + m 3)Kh - *»s)! "] ] ( 71 +72 + 73 + 1)!( Jl -72 + /s)Kii + 72 -h)K ~/l +7*2 - "* 3 )!( j 3 + m 3 )! J Vi -7i- ( 2 7i) ! ( -7*1 +7*2 +73)!(ii +72 + ^3)!(j3 - «s)! "1 1/2 . (jl +72 +73 + l)!(j'i -72 +73)Kii +7*2 -ja)K -jl +h ~ ™3)KJ3 + «b)L are obtained directly from (106.14). The derivation of the formula ' (ji +72 -J3)K h-h+J3) K -j\ +72 +73)! ~| 1/2 (2?+ 1)1 /7i 72 73\ = r Qi+72-73)!0i-72+73)!(-7i+72+73)H ] \0 0/ L (2/>+l)! J p\ - , (106.18) (p-h)KP-h)KP-33)\ where 2p = j\ +72 +73 is even, requires a number of additional calculations ;f when 2p is odd, this 3/-symbol is zero owing to the symmetry property (106.6). Table 9 gives for reference the values of the 3/-symbols for 73 = ^, 1, § , 2. For each js the minimum number of 3/-symbols are shown from which the remainder may be obtained by means of the relations (106.5), (106.6). Table 9 Formulae for 3j -symbols (' +i '' *\ = (-nn-1/.r J '- m+i T \ m -m-i i/ L(2/+1)(2;+2)J (-D-f '' ') \m —m—ms W3/ t See Edmonds' book already quoted. §106 3j-symbols 407 \«8 n V«3 y+i ^3 n j+ j+4 2m [2/(2/+l)(2/+2)]i/2 2(y+m+l)(y-m+l)-ji/ 2 r 2Q+m+ix;-m+in ;+ L(2/-+l)(2/-+2)(2y+3) J F(/-m)(y+m+l)-| 1/2 2y(2y+i)(2y+2) J _ r U-™)U- m + 1 ) 1 1/2 L(2y+i)(2y+2)(2y+3)J (-ly-m+l^p 1 J ? \ \m —m—ms m%) r /— m+h ~| 1/2 y+i -(y+3m+f) 3 2 U 2 l2/(2/+l(2y+2)(2y+3)J 2y(2y+i)(2y+2)(2y+3). r 3(y-m+ixy-m+f)(y+m+§) -ji /2 ' ' L (2y+i)(2y+2)(2y+3)(2y+4) J ma I _ r 3(y-m-|)(y-m+|)(y+m+|) -|i /2 L 2y(2y+i)(2y+2)(2y+3) J . a r (y-m-i)0-m+|)(y-m+f) -ji/2 ;+ " L (2y+i)(2y+2)(2y+3)(2y+4) J (-D-f j 2 ) \W3 11 . 2[3m2-y(y+l)] [(2y- 1)27(2; + l)(2y+2)(2y+ 3)F 2 ;+ m L2y(2y+ l)(2/+2)(2/ + 3)(2/+4)J r6(y+m+2)(y+w+l)(y-m+2)(y-w+l)-] 1 / 2 3+2 L (2y + 1)(2; + 2)(2j+ 3)(2y+ 4)(2/+ 5) J 408 Addition of Angular Momenta §107 ^^3 1 1" 6(j+m+l)(j-m) -ji /2 m) l{2j- l)2j(2j+ l)(2/+2)(2/+3)J n -2 ( y + 2, + 2)r (i-^DO--^) _1 1/2 12/(2/+ l)(2/+2)(2/+3)(2/+4)J 3 L (2y+l)(2/+2)(2y+3)(2y+4)(2y+5) J \w 3 2 ;i i+i 6(y-w-i)(y-w)(y+m+i)(y+m+2)-i 1 / 2 (2;-l)2/(2y+l)(2; + 2)(2; + 3) J ■(y-m-l)(y-m)(y-m+l)(y+m+2)"] 1 / 2 2y(2y+i)(2y+2)(2y+3)(2y+4) [ L 2y(2y+i)(2y+2)(2y+3)(2y+4) J • 2 r (y-m-l)(y-m)(y-m+l)(y-m+2) ni/2 ;+ L (2/+l)(2; + 2)(2;+3)(2/+4)(2/+5) J PROBLEM Determine the angle dependence of the wave functions of a particle with spin J in states with given values of the orbital angular momentum /, the total angular momentum j and component thereof m. Solution. The problem is solved by the general formula (106.8), in which ^'must be taken as the eigenfunctions of the orbital angular momentum (i.e. the spherical harmonic functions Yim t ), and t/r< 2 > as the spin wave function x( a ) (where a = ±£): y, w = (-1)m^i/V(2;+1)T( l * 3 )Yi, m - (T x(<y). *-~ ' \m—a a —ml <r Substituting the values of the 3/-symbols, we obtain 7j+m ij—tn -TT-xCi) Yi, m -U2 + J —rx(-i) Yl,m+l/2, lj—m+1 fj+m+1 *i- 1/a> „ = _^_— ^)y, >w _ 1/2+ ^__ x (- i) y ZiWl+1/2 . §107. Matrix elements of tensors In §29 formulae have been obtained which give the matrix elements of a vector physical quantity in terms of the value of the angular momentum component. These formulae are really a particular case of the corresponding general formulae for an irreducible (see §58) tensor of any rank.f t The analysis of the problems discussed in §§107-109, and most of the results given, are due to G. Racah (1942-1943). §107 Matrix elements of tensors 409 The set of 2k + 1 components of an irreducible tensor of rank k (an integer) are equivalent, as regards their transformation properties, to a set of 2k +\ spherical harmonic functions Y* a , q = —k,...,k (see the last footnote to §58). This means that, by means of appropriate linear combinations of the components of the tensor, we can obtain a set of quantities which are trans- formed under rotations as the functions Y k q. A set of such quantities, which will be denoted here byfjcq, is called a spherical tensor of rank k. For example, k = 1 for a vector,