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TEXT BOOK
THERMODYNAMICS
BY
PAUL S. EPSJEIN
PROFESSOR OF THEORETICAL PHYSICS, CALIFORNIA INSTITUTE Of
TECHNOLOGY
NEW YORK
JOHN WILEY VSONS, INC.
LONDON: CHAPMAN 84 HALL, LIMITED
COPYRIGHT, 1937, BY
PAUL S. EPSTEIN
All Rights Reserved
This book or any part thereof must not
be reproduced in any form without
the written permission of the publisher.
PRINTED IN U. 8. A.
="=*ES8 OF
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TEXTBOOK OF THERMODYNAMICS
CHAPTER I
THE THERMAL PROPERTIES OF MATTER
1. The concept of temperature. Heat and cold belong to the
most fundamental of our sensations. From time immemorial the
phenomena connected with them received a great deal of practical
attention because of their vast importance in everyday life, but they
were not made amenable to scientific treatment until a comparatively
recent date. The possibility of doing this is due to the following
empirical facts: (1) When a hot and a cold body are brought in con
tact, the former gets cooler, the latter warmer. After the lapse of a
sufficiently long time a stationary state is reached in which no further
changes occur. We say, then, that the two bodies are in thermal
equilibrium. This equilibrium persists if the bodies are separated
and then brought together a second time (provided they were not
exposed, in the meantime, to the influence of other hotter or colder
bodies). (2) Two bodies which are separately in thermal equi
librium with the same third are in equilibrium with each other. We
say, then, that they have the same temperature.
The second fact permits us to use one suitable body as a stand
ard or "thermometer" with which to compare the temperatures of
others. In order to determine the degree of heat of the thermometer,
one, usually, has recourse to the property of matter of changing its
size or volume when heated. In this way a scale of temperatures is
established. The current types of thermometers are well known and
need not be described here.
When brought into equilibrium with any body, the thermometer
associates the degree of heat of this body with a definite reading of its
scale. It replaces the vague notion of hot or cold by a precise numerical
datum. Its invention was, therefore, the necessary antecedent of
a scientific analysis of the phenomena of heat. (Compare
section 39).
2 TEXTBOOK OF THERMODYNAMICS 1 3
2. Subject of thermodynamics. Definitions. Thermodynamics
deals with systems which, in addition to mechanical and electromag
netic parameters, are described by a specifically thermal one, namely,
the temperature or some equivalent of it. We have seen that for the
very definition of temperature one must invoke the concept of thermal
equilibrium. Thermodynamics is, therefore, of necessity, essentially
a science about the conditions of equilibrium of systems and about
the processes which can go on in states little different from the
state of equilibrium. We introduce the following terms which, for
the most part, are those in general usage:
A homogeneous system is uniform in every part, with respect both to
its chemical composition and to its physical condition. The fact men
tioned in the preceding section that two bodies set themselves into
thermal equilibrium applies, of course, also to two adjacent parts of
the same body. The statement that the temperature of a homogeneous
system is uniform throughout is, therefore, equivalent to saying that
it is in equilibrium.
A simple homogeneous system or, for short, a simple system is one
completely defined by the three parameters V (volume), p (pressure),
and T (temperature). Such are gases and liquids in a static condition
and solids in a state of isotropic stress removed from the action of
gravitational and electromagnetic forces. The case of one or several
simple systems is, by far, the most important. Simple systems offer
the most interesting applications, and their treatment is sufficiently
typical to bring out all the methods of thermodynamics so that a
generalization is easily made when necessary. We shall, therefore,
restrict ourselves in many of our discussions to simple systems, and
shall take up the more general case only when this is required for
special applications.
A heterogeneous system is composed of a number of homogeneous
ones separated either by surfaces of discontinuity or by wall partitions.
Those homogeneous parts are called the phases of the system.
We shall admit the existence and the use of partitions of different
types of which the most important are the heatconducting one which
permits two bodies separated by it to set themselves into thermal
equilibrium, and the insulating or adiabatic partition which prevents
any exchange of heat.
3. The thermal equation of state. We have mentioned in section 1
that the temperature of a thermometer is usually inferred from its
extension. The deeper reason why it is possible to do so lies in a very
fundamental and general property of every simple system: there
exists a functional relation between the temperature, the volume, and
I 3 THE THERMAL PROPERTIES OF MATTER 3
the pressure which is called the equation of state, or more accurately,
the thermal equation of state. This relation between T, , and 7
can be written in the form
pf(V,T). (1.01)
Each of the three variables can, therefore, be regarded as a function
of the other two. Subjecting the variables to small changes, we can
form the following three partial derivatives:
(1) We keep the pressure constant (A/> = 0) and give to 7 and T
the small increments A 7 and AT. The limit of the ratio, for AT 1 = 0,
is the partial derivative lim(A7/Ar) = (97/37%. (It is customary
to indicate the parameter which is kept constant as a subscript.)
Divided by the volume 7, this partial defines the coefficient of thermal
expansion
< /^TA
(1.02)
which represents the relative (i.e. referred to unit volume) increase of
volume per unit increase of temperature, at constant pressure.
(2) Keeping the temperature constant (AT" = 0), we obtain in a
similar way the partial (QV/Qp)T connected with the coefficient of com
pressibility
1 /^T/\
(1.03)
which gives the relative decrease of volume per unit increase of
pressure, at constant temperature.
(3) When the volume remains constant (A 7 = 0), the partial
(dp/dT)v is obtained, related to the coefficient of tension
(1.04)
The inverse partials are, simply, the reciprocals, (37/97%
= l/(9r/97)p, etc., and do not give us anything new, and even the
three partial derivatives which we have formed are not independent
as there exists a relation between them. This becomes apparent
when we write down the general expression for the total infinitesimal
change A/> which results when both the volume and the temperature
are changed by infinitesimal amounts. It is, according to the rules of
differentiation,
'3P\ ,rr , (&
4 TEXTBOOK OF THERMODYNAMICS 1 4
In the special case, A = 0, the ratio AF/AT 1 is, precisely, the
partial (dV/dT) p . Since the left side of the equation vanishes in
this case, we can solve it with respect to this ratio and find
This relation can be regarded as a differential form of the equation
of state (1.01).
In the general case of a homogeneous system described by other
parameters (in addition to temperature and volume) there exist
several equations of state as will become clear in section 7.
Exercise 1. In nitrogen at t = 20 C and p = 0.134 atm, a = 0.00368 and
ft = 0.00368. Calculate the compressibility 0i.
Exercise 2. Suppose that one of the equations of state of a homogeneous
system depending on many variables has the form
The increment of y, analogous to the expression for Ap, is
Show that by keeping all the variables y, Xi, . . . X n , T constant, except two, one
can obtain n(n f l)/2 new relations of the type (1.05) between triples of partials.
(Remark: The set of relations so obtained includes all that are independent. From
them there can be derived a great number of other (dependent) relations of a
similar form.)
4. The perfect gas. Absolute temperature. The socalled perma
nent gases of nature follow, more or less closely, a number of well
known laws. The first of these is the BoyleMariotte law according
to which, for a constant temperature, the product of pressure and
volume remains constant
pV  const. (1.06)
When we inquire about the behavior of a gas at varying tempera
tures, it is convenient to use a gas thermometer filled with this same
gas as the thermometric fluid. The pressure of the thermometer is
kept constant (o), and the zero of the temperature scale (t 0)
is chosen arbitrarily, say, as that corresponding to a volume Fo.
I 4 THE THERMAL PROPERTIES OF MATTER 5
The temperature t is, then, defined as proportional to the relative
increase of the volume over Fo,
or
V  7 (1 + <**)* (1.08)
a being a constant.
According to the law of CharlesGayLussac, the formula (1.07)
is universal for all permanent gases: Used as thermometric fluids in
the same thermometer, they give, within a good approximation, the
same readings, or, in other words, they have all pretty closely the same
coefficient of expansion a. The uniformity can be best judged in the
following way. Suppose we use a centigrade scale: that is, we choose
the temperature of melting ice as / = and that of boiling water as
t = 100. Equation (1.08) gives, then, for the volume at t = 100,
Fioo = F (l + lOOa), whence a = (Fioo  Fo)/100F . The quan
tity defined in this way is called the mean coefficient of expansion,
between and 100, and is denoted by o,ioo. In general this quan
tity depends on the pressure po kept constant during the expansion.
The mean coefficients of expansion <*o, 100 were investigated by Henning
and Heuse who measured them with considerable accuracy in helium,
hydrogen and oxygen. These measurements are well represented by
the expressions of Table I, 1 where po is expressed in atmospheres.
Degrees of the centigrade scale will be indicated by the symbol C.
TABLE 1
Gas <>, too X 10 7
He ......... 36604  25p
H 2 ......... 36604  16p Q
N 2 ......... 36604 + 167/>
We arrive at a more rational choice of the zero point of our scale
if we displace it by I/a, defining the new temperature as
T = t + I/a. (1.09)
The position of the old zero point / = on the new scale is then
Jo = I/a, and eq. (1.08) takes the form V TtPo F r/7o. The
temperature T is proportional to the total volume of the gas. We
substitute this into eq. (1.06) whidi we refer to the temperature T 9
* E. Kenning and W. Heuse, Zs. Physik 5, p. 285, 1921.
6 TEXTBOOK OF THERMODYNAMICS 1 4
writing pV = poV TtPo , and obtain a formula representing the com
bination of the BoyleMariotte and CharlesGayLussac laws:
pV CT, (1.10)
where C is an abbreviation for the constant C
A simple system in equilibrium is completely homogeneous and
uniform in all its parts (compare section 2). The properties of our gas,
in every small region of it, are, therefore, completely defined by the
temperature T and pressure p and cannot depend on its total exten
sion or mass. If we compare two specimens of the same gas of dif
ferent masses, but at the same temperature and pressure, their respec
tive volumes V must be, obviously, proportional to their masses. It
follows then from eq. (1.10) that the constant C must be also propor
tional to the mass of the gaseous system under consideration.
We shall use as the unit of mass of a chemically homogeneous or
pure substance the mol, also called the grammolecule. If /* denotes
the molecular weight of this substance, its mol contains just n grams.
To simplify our terminology we shall apply the term " molecular
weight " also to monatomic substances where it has the same meaning
as " atomic weight/' Let us consider a homogeneous system which
consists of N mols of a pure substance. The volume occupied by one
mol of it is called the specific molal volume or, simply, the molal volume
and is denoted by v.
We can apply this to our gas system if it is chemically homogeneous.
The total volume V and the constant C can be represented as
V**Nv, C = NR, (1.11)
where R is an abbreviation for R = poVo/To or the constant C referred
to one mol of the gas. The equation of state (1.10) takes then the form
pV = NRT, (1.12)
or
pv = RT. (1.13)
The third empirical law which we have to invoke is that of Avo
gadro. It tells us that the molal volume v is, within a close approxima
tion, the same for all permanent gases, namely, under normal condi
tions (p = 1 normal atm = 1.013249 X 10 6 dyne cm 2 , t = C)
it is vo = 22414 cm 3 . Consequently, the gas constant R is also uni
versal. It is an immediate consequence of the definition of the mol,
that 1 mol of any substance contains the same number n A of molecules,
which is called the Avogadro number. It has been determined with
I 4 THE THERMAL PROPERTIES OF MATTER 7
fair accuracy and has been found to be n A =* (6.064 =fc 0.006)
X 10 23 mol" 1 . A muchused quantity is the ratio
*  R/n A9 (1.14)
which represents the gas constant per molecule and is called the Boltz
mann constant. With its help eq. (1.13) can be written in an alterna
tive form. If we denote by Z the total number of molecules in N mols
of a gas having the volume F, and by z = Z/F the number per unit
volume, we can write Z = Nn A , NR = Nn A k = Zk = Vzk, and
from (1.12)
p = zkT. (1.15)
If, instead of a single gas, we consider a mixture, we must again
have recourse to experience. Observations on mixtures of permanent
gases show that, with a good approximation, each of the constituent
gases behaves as if the others were not present. In the case of <r
gases contained in a vessel of the volume Fin the respective quantities
of Ni, . . . N 9 mols, each of them exerts a partial pressure upon the
walls of the vessel which can be computed in the same way as if the
gas were single, i.e. by the formula
p h V = N h RT. (1.16)
The total pressure of the mixture is the sum of these partial
pressures
P = pi + p2 + ...+p., (1.17)
and its equation of state is obtained by summing the formulas (1.16)
for h = 1, 2, ... <r
(Ni + ... + N ff )RT. (1.18)
None of the above laws is followed by any of the permanent gases
quite rigorously, and the formula (1.18) which embodies all four of
them is, therefore, only an approximate expression of the facts. If
we compare two gas thermometers filled with different gases and
having the same fixed points, the readings between the fixed points
will not be strictly identical because the rate of expansion is, in reality,
not constant but slightly dependent upon the temperature, in a dif
ferent way for the different gases. The same thing is true for non
gaseous thermometric fluids like mercury and alcohol. The scale of
temperatures established by any particular thermometer construction
is, therefore, special and arbitrary. In order to have a scale which
does not depend on the individual properties of any special substance,
we imagine the socalled perfect gas: an ideal fluid which strictly
8 TEXTBOOK OF THERMODYNAMICS 1 4
follows the equation of state given by formula (1.12). Let us, further,
imagine a thermometer filled with a perfect gas and let us specify its
scale in such a manner that between the points of freezing and of
boiling water (both at 1 norm atm = 760 mm Hg) are 100 equal
divisions. The temperature which we would measure with such a
thermometer is called the absolute temperature. We shall, consistently,
denote it by T (capital), reserving / (small) for the temperature mea
sured in an arbitrary scale.
It might be thought, at first sight, that such a definition of the
absolute temperature must be futile since the perfect gas, and the
perfect thermometer filled with it, are ideals which do not exist in
nature. We shall show, however, in section 29 that there is a simple
procedure to determine the absolute temperature, in every case, by
indirect measurements. Every thermometer can, therefore, be cali
brated in the absolute scale within the accuracy with which it can be
read, and such calibrated thermometers are, in fact, readily available.
Making use of the results of this calibration, we can state that the
point of freezing water corresponds to TQ = 273.1 of the absolute
scale. Combined with the value of the molal volume mentioned above
(VQ 22414 cm 3 ), this leads to the following numerical values for the
gas constant: R = (82.049 0.009) atm cm 3 deg" 1 mol^ 1 (if the
pressure is measured in normal atmospheres) or R = (8.3136 0.001)
X 10 7 erg deg" 1 mol" 1 ) (if it is measured in dynes per cm 2 ). For
Boltzmann's constant defined by (1.14) there follows the numerical
value k = 1.371 X 10~ 16 erg deg" 1 . The question how closely the
permanent gases approximate the behavior of a perfect gas will be
more fully discussed in section 6. Here we shall only state that,
according to their degree of perfection, they can be arranged in the
following order: carbon dioxide, oxygen, air, nitrogen, argon, hydro
gen, helium. Carbon dioxide is the least and helium the most nearly
perfect gas of this series. The deviations of helium from eq. (1.13) of
perfect gases are, in fact, quite small under most conditions. Except
at extremely low temperatures, helium can be used as the perfect
thermometric fluid unless a very high accuracy is required. (Compare
also section 29.)
Exercise 3. Show that the coefficients of thermal expansion and of compres
sibility are for a perfect gas
I 5 THE THERMAL PROPERTIES OF MATTER 9
and check the validity of the relation (1.05). Calculate the coefficients numerically
(per 1 deg, and 1 mm Hg as units) at normal conditions: r 273M, p
760 mm Hg.
5. The equation of Van der Waals. The next step in the approxi
mation to the properties of real matter is represented by an equation
given by Van der Waals. In spite of its simplicity, it comprehends
both the gaseous and the liquid state and brings out, in a most remark
able way, all the phenomena pertaining to the continuity of these two
states. This equation has the form
where a and b are two numerically small constants and v is the molal
volume. It was derived by statistical considerations of so simple a
nature that, even in a book devoted to thermodynamics, we can say a
few words about the interpretation of the constants. It is shown in
the kinetic theory that eq. (1.13) corresponds to a gas whose molecules
are material points which do not collide or otherwise interact with one
another. If we consider a gas with molecules of finite size, mutually
exclusive as to their extension but not interacting in any other way,
the difference is that the centers of the molecules cannot spread out
in the whole volume of the gas but only in that part of it which is
not occupied by other molecules and not immediately adjacent to
them. In the first approximation, for the molal volume v there must
be substituted the covolume, v b, where the constant b is proportional
to the sum of the volumes of all molecules in one mol of the gas. The
equation of state becomes, therefore, p = RT/(v b). On the other
hand, if the molecules of the gas do interact at a distance, say, attract
one another, there must be added to the external pressure p the
internal pressure due to this attraction. Whatever its law is, it is
safe to assume that two volume elements of the gas ATI and Ar2
act upon each other with a force Fi2 proportional to the product of
the masses in them, i.e. Fi% cc AwiAma. If we choose the mol as
the unit of mass, we have Ami = ATI/IF, Aw2 = Ar2/v, so that
F\2 cc ATiAT2/u 2 . This means that, when the density of the gas in a
given vessel is changed by adding more gas or subtracting it, all the
internal forces change in the ratio 1/v 2 . Since the pressure is defined
as the force per unit area, this applies also to the internal pressure
and we obtain for it the expression a/v 2 which is added to p in Van
der Waals' equation. 1
1 It must be assumed, also, that the force Fit decreases rapidly with the cfistance
between the volume elements An and An. Otherwise the coefficient a would not
10
TEXTBOOK OF THERMODYNAMICS.
15
Let us plot the pressure p against the molal volume v keeping the
temperature T' in eq. (1.21) constant. Such a curve is called an
" isothermal." As long as the molal volume v is very large, b is negli
gible compared with v, and a/v 2 compared with p. In the region of
large volumes the curve is, therefore, identical with that given by
the equation of perfect gases, namely, a hyperbola asymptotically
approaching the horizontal axis p = for v = o . When the volume
is very small and close to its lower limit 6, the second term on the
right side is again negligible compared with the first and the curve
becomes again hyperbolic. However, its asymptote is not the vertical
V = (as in the case of the perfect gas) but the vertical v = b. The
shape of the isothermals in the intermediate range is given in Fig. 1.
They are curves with one minimum
and one maximum which are real for
low temperatures and complex for high
so that, above a certain temperature
T c (critical temperature), the pressure
is, in them, a monotonically decreas
ing function of the volume (dp/dv <[0).
Below T c the curves are not con
tinually decreasing but have a middle
part (between the minimum and the
maximum) in which p increases with
v, (dp/dv > 0). It is obvious that this
middle part, dotted in our curves,
can have no physical reality. In fact,
let us imagine the fluid in a state cor
responding to this part of the curve
contained in a heatconducting vertical cylindrical vessel whose top is
formed by a piston. The piston can slide up and down in the cylinder,
and we put on it a load exactly balancing the pressure of the gas.
If we take a little weight off the piston, there will no longer be
equilibrium and it will begin to move upward. However, as it moves
the volume of the gas increases and with it its pressure. The resultant
force upon the piston gets larger, retaining its upward direction. The
piston will, therefore, continue to move and the gas to expand until it
reaches the state represented by the maximum of the isothermal. Vice
versa, if we add ever so little to the load of the balanced piston, the
gas will collapse to the state corresponding to the minimum of the
be an internal constant of the gas but would depend on the shape and size of the
container. It does not follow from the above argument and is a separate assump
tion that a is independent of T and p.
o
FIG.
1. Isothermals according
Van der Waals 1 theory
v
to
I 5 THE THERMAL PROPERTIES OF MATTER 11
isothermal. The relation dp/dv g is, therefore, the condition of
dynamical stability, and in regions where it is not satisfied the fluid
is in an absolutely unstable state.
Leaving out the physically unreal dotted portion, the low tempera
ture curves break up into two disconnected branches. The one of
small molal volume corresponds to the liquid state of matter while
that of high volume represents its gaseous state. The important fact
that the same equation applies to both states of aggregation was
called by Van der Waals the continuity of the liquid and gaseous states.
Experiments show that the following things actually happen when we
gradually increase the volume (in a heat bath of constant temperature),
starting from a point A of the liquid state (Fig. 1). The expansion
goes along the liquid branch of the isothermal up to a certain point B.
In this state the liquid begins to evaporate so that the system is no
longer homogeneous, but heterogeneous, consisting of a liquid and a
gaseous phase. As the volume is further increased, more and more
liquid is evaporated while the specific properties of the two phases
remain unchanged. In this part of the isothermal expansion the
pressure (called the boiling pressure at the temperature 7') is constant,
so that it is represented by a horizontal line, and remains so until all
the liquid is evaporated (point E). From then on the system is again
homogeneous and expands along the gas branch EF of the isothermal.
Why the system behaves in such a way, and where the point B lies,
are typical problems of thermodynamics which will be answered at
length in section 43. The part BC of the liquid branch (like DE of
the gas branch) represents only relatively unstable states: the vapori
zation of a liquid is greatly facilitated by the presence in it (or on its
surface at the walls) of little bubbles of foreign gases. If these gases
are driven out by a preliminary thorough boiling of the liquid and a
suitable heat treatment of the vessel, it is possible to make the system
expand along the liquid branch beyond the point B, bringing it into
the socalled superheated state. In this state the pressure may become
negative, in other words, the liquid is able to support a certain
amount of tension; but at the slightest provocation it will jump
over into the thermodynamically stable state on the straight
line BE.
Above the critical temperature there is no discontinuity in the iso
thermal, and it is arbitrary whether to call the fluid a liquid or a gas.
Usually, it is considered a gas and, then, the statement applies: // is
impossible to liquefy a substance at temperatures higher than the critical.
If the substance follows eq. (1.21) of Van der Waals, its critical tem
perature can be determined in the following way. Below it, the iso
12 TEXTBOOK OF THERMODYNAMICS I 5
thermals have a maximum and a minimum which are analytically
determined by the condition
+ 2_* o. (1.22)
As the temperature rises, the maximum and the minimum draw
closer and closer together, and, in the critical isothermal, they coincide,
forming a point of inflection. This point corresponds to the socalled
critical state of the substance characterized by the values p c , v c , T c of
pressure, molal volume, and temperature which are known as the
critical constants. The analytical condition for a point of inflection is
_ 6a
^ )
In the critical point all three equations (1.21), (1.22), (1.23) must
be satisfied. It is, therefore, easy to express the critical constants in
terms of the Van der Waals constants a, b, R by solving these three
equations with respect to the three variables p, v, T. The result is
..
One can use the critical constants as the units in which to express,
respectively, the pressure, specific volume, and temperature. The
state of the substance is then described by the socalled reduced variables
p v T .
ft * = :: T = V (L25)
pc V c lc
If we substitute these quantities for p, v, T and, at the same time,
replace the Van der Waals constants by the critical, the equation of
state (1.21) takes the form
Referred to the reduced variables, the equation does not contain
any specific constants and is, therefore, the same for all Van der Waals
substances. States of two substances in which they have the same
respective values of T, <f>, and T are called corresponding states, and
eq. (1.26) expresses the law of corresponding states: If two of the
reduced variables, IT, ^>, T, are the same, respectively, for different sub
stances the third is also the same and their states are corresponding. The
mathematical reason for the validity of the law of corresponding
I 6 THE THERMAL PROPERTIES OF MATTER 13
states in connection with Van der Waals' equation is that it contains
just as many constants as variables, namely three. However, the
law of corresponding states is not restricted to Van der Waals sub
stances. It is often satisfied with a fair degree of precision in groups
of substances which show marked deviations from eq. (1.21), as will
be shown more fully in the next section.
It is well to add here a few words about gas mixtures. The additiv
ity of partial pressures which exists in perfect gases, according to
eq. (1.17) does not hold in mixtures of Van der Waals gases. The
physical reason for that additivity is that each perfect gas behaves as
if the others were not present, and this means that there is no inter
action between their molecules. On the contrary, the aim of Van der
Waals' theory is, precisely, to take into account the interaction.
For a binary mixture (i.e. consisting of two gases) the kinetic theory
has been carried through by Lorentz. 1 The equation of state obtained
by him for such a mixture is formally identical with the ordinary
Van der Waals eq. (1.21), but the symbols v t a, b have a different
meaning. If the mixture contains NI mols of the first gas and N 2 of
the second, v = V/(N\ + N 2 ), while the quantities a and b depend
on the mol fractions Xi = N\/(N\ + N 2 ) and x 2 = N 2 /(Ni + N 2 )
and have the form
a = #n xi 2 + 2ai2 Xix 2 f 022 x$
b = bn xi 2 + 2bi2 xix 2 + b 22
(1.27)
The coefficients an #22 are those of the pure gases, while #12
characterizes the mutual action ; the same applies to the coefficients b.
Exercise 4. Calculate by means of eqs. (1.24) the Van der Waals constants
a, b and the ratio RT c /p c v c for the following substances from the critical data
(R 82.07):
Methyl acetate 506.9 46.33 227.6
Methyl formate 487.2 59.25 172.0
Exercise 5. Derive the coefficients of thermal expansion (1.02) and of compres
sibility (1.03) for the equation of Van der Waals. (Remark: as this equation is
readily resolved with respect to />, it is convenient first to express a and pi in terms
of partials of p t by means of (1.05) or in any other way.)
6. Behavior of real substances. As far as gases are concerned, it
was pointed out in the preceding section that eq. (1.13) of a perfect
1 H. A. Lorentz, Ann. Physik 12, p. 127, 1881. For a discussion of this equation,
in relation to experimental data, see: J. D. Van der Waals, Continuity of the
Gaseous and Liquid States. Vol. II. German edition of 1900.
14
TEXTBOOK OF THERMODYNAMICS
I 6
gas is already a good approximation as long as the molal volume v is
large. Although the equation of Van der Waals has its limitations
(see below), the numerical values of its constants a and b are apt to give
an idea how much a given substance deviates from the law of perfect
gases and indicate the order of accuracy here involved. We give in
Table 2 these constants and the critical data of a number of gases.
STABLE 2
Gas
Chemical
Symbol
t c *C
PC (atm)
b = \v c
(cm')
a X 10
(atm cm 6 )
Helium
He
267.9
2.25
23.71
0.03415
Neon
Ne
228.35
26.85
17.10
0.2120
Hydrogen
H 2
239.9
12.8
26.61
0.2446
Argon
A
122.4
48.0
30.22
1.301
Nitrogen
N 2
147.13
33.5
38.52
1.346
Oxygen
O 2
118.82
49.7
32.58
1.361
Carbon monoxide
Carbon dioxide
CO
CO 2
138.7
31.0
34.6
72.9
39.87
42.69
1.486
3.959
Nitrous oxide
N 2 O
36.5
77.66
44.18
3.788
Water vapor
H 2 O
374
217.5
30.52
5.468
Chlorine
C1 2
144
76.1
56.26
6.501
Sulfur dioxide
SO 2
157.15
77.65
56.39
6.707
The constant ft, depending on the molecular volume, is not greatly
different for the different gases. The relative correction in the magni
tude of the pressure, due to the presence of this constant in eq. (1.21),
is roughly b/v. In the neighborhood of the normal conditions
(/ = C, p = 1 atm), the molal volume is about 22 400 cm 3 /mol
and the correction has the numerical value of 0.001 to 0.002. On the
other hand, there is a considerable variation, from gas to gas, in the
constant a, and the correction a/v 2 depending on it varies widely.
For helium under normal conditions it is only 0.00007, for nitrogen
and oxygen 0.0027. For the latter two gases, it is of the same order
of magnitude as the first correction (0.0017 and 0.0015), and, since
its sign is opposite, the total correction is reduced. For water steam
at the boiling point (t = 100 C, p = 1 atm), it becomes 0.006.
It was known already to Van der Waals himself that his equation
does not give an accurate representation of the properties of a fluid
in all ranges of temperatures. In fact, it is possible to determine the
coefficients a and b from the data about pressures and volumes relating
to a single isothermal. When the coefficients, obtained in this way
for widely different isothermals, are compared, they turn out not to
16
THE THERMAL PROPERTIES OF MATTER
15
be the same, especially, a is found to increase as the temperature is
lowered. Another indication that eq. (1.21) is not accurate in the
vicinity of the critical state is as follows. Theoretically there exists
the following relation between the critical data p e , v c , T if the expres
sions (1.24) are valid:
K
RT C 8
  2.67.
(1.28)
In reality, the ratio of the empirically measured critical data
RT c /pcV c is nearly the same within large groups of substances but it
has not the value 2.67 required by the theory. The constancy of the
ratio R is an indication that the law of corresponding states still
approximately holds when the Van der Waals equation fails. In
Table 3 we give the critical data of a few organic substances.
TABLE 3
Substance
Formula
To
PC
v e
RT c /p c v
Ethyl ether
C 4 Hi O
467.0
35.6
281.9
3.814
Benzene
CH
561.7
47.89
256.2
3.755
Bromobenzene
CflH 6 Br
670.2
44.64
323.4
3.813
Chlorobenzene
C 6 H 6 C1
632.4
44.62
307.9
3.776
Ethyl formate
CaHeO,
508.5
46.76
229.0
3.895
Ethvl acetate
C 8 H 8 O2
523.3
38.00
286.0
3.885
nPentane
C*5l~l 1J
470.4
33.04
310.9
3.762
tPentane
v^5i~i 12
461.0
32.92
307.3
3.734
Octane
CaHig
569.4
24.65
490.9
3.864
Heptane
C 7 Hi
539.4
26.88
537.8
3.945
Hexane ....
(.gH 14
507.4
29.62
366.9
3.830
Carbon tetrachloride ....
C C1 4
556.3
44.98
275.9
3.677
Tin tetrachloride
SnCl 4
591.9
36.95
351.2
3.755
In spite of the shortcomings just mentioned, the Van der Waals
equation is extremely useful. It gives an excellent quantitative account
of the behavior of gases whose density is not too high. Beyond this
it offers a qualitative picture of the phenomena of condensation,
critical state, etc. However, it must be always remembered that it
cannot be used for quantitative predictions of phenomena involving
high densities. Many other equations of state were proposed but,
in part, they apply only to a narrow range of the variables, in part,
they offer a gain in accuracy insufficient to compensate for the much
more complicated form. Therefore, we shall mention here only two
of them.
16 TEXTBOOK OF THERMODYNAMICS I 6
(A) The mean empirical, reduced equation of Kamerlingh Onnes 1
is expressed in terms of the reduced variables (1.25),
K is the ratio (1.28) of the critical data, while the coefficients
depend on r
iii #2, . . being constants. The other coefficients have the same
structure.
This equation grew out of the studies of its author on the validity
of the law of corresponding states. He found that there exist large
groups of substances which obey the law of correspondence to the
following extent. The substances of a group can be arranged in a
series so that each of them has a close correspondence with the pre
ceding substance while the series, as a whole, shows progressive and
considerable changes in the constants. The above equation applies
to a fictitious substance representing the mean type of the series. The
results of Kamerlingh Onnes show us two things: (1) how extremely
complicated the conditions become when a high accuracy is aimed at;
(2) that the law of corresponding states has an approximate validity
beyond that of any of the simpler equations of state.
(B) The equation of state of Beattie and Bridgeman 2
A
* v 2 ^ ' ~ l v 2 '
[ e = c/vT*,
AQ, Bo, a, i, c being constants. This equation is one of the most recent
that have been proposed so that its authors had the benefit of the
experience of their predecessors and of more accurate experimental
data. 3 It gives good results over a wide range of the variables.
1 H. Kamerlingh Onnes and W. H. Keesom, Die Zeistandsgleichung. Encyklo
paedie math. Wiss. Vol. VI, Article 10.
2 J. A. Beattie and O. C. Bridgeman, Proc. Am. Acad. 63, p. 229, 1928.
1 Another of the more recent equations is that due to A. Wohl (Zs. physik.
Chemie 87, 1, 1914; 99, p. 207, 1921). It is often quoted, but the writer has diffi
culty in understanding it, as he is not convinced that it satisfies the requirements of
an equation of state.
tl  e) ,. . A B _ ^ _
16
THE THERMAL PROPERTIES OF MATTER
17
The above equations of state apply, primarily, to gases, secondarily,
to liquids at temperatures not far from the critical. It is not necessary
to say much about matter in the condensed state. The thermal
expansion and compressibility
of condensed substances are
so small (compared with
those of gases) that they are
not important in most ther
modynamical applications.
In the first approximation,
they are altogether neglected ;
in the second, they are taken
as constant over finite ranges.
The process of transition from
the liquid to the gaseous state
is in real substances the
same as that described in
connection with the Van der
Waals equation and illus
trated by Fig. 1. As an
example we give in Fig. 2 the
actually observed isothermals
in carbon dioxide. (The mp FlG 2 .Empirical isothermals in carbon
lal volume is given here in
fractions of the normal vol
ume at C and 1 atm, v,
marked by the symbol X).
30
D .005 ".01.
V/Vn
FIG. 2. Empirical isothermals
dioxide.
= 22 256 cm 3 ; the critical point is
Exercise 6. Calculate the critical data in terms of the constants of the equation
of state
/ n \
 b) RT (Berthelot)
by way of the formulas (1.22), (1.23), (1.24).
Exercise 7. Do the same for the equation
p(v b) = RT exp (c/RTv). (Dieterici)
(Remark: when the first and second partials of p vanish, those of log p also vanish.)
CHAPTER II
THE FIRST LAW OF THERMODYNAMICS
7. The element of work. The concept of work is defined in
mechanics as follows. When a material point moves under the action
of a force F, through the infinitesimal distance dl in the direction of F,
the elementary work done by this force is Fdl. When the directions
of the path dl and of the force do not coincide but include an angle a
only the projection of F on the direction of motion Fi = F cos a con
tributes to the work, and the element of work has the expression
DW = Fidl = Fdl cos a = F x dx + F y dy + F,dz, (2.01)
if we denote, respectively, by F x , F v , F, and dx, dy, dz the projections
of F and of dl upon cartesian axes of coordinates. In the case of the
general dynamical system of n degrees of freedom, the element of
work is expressed in a similar way
DW  yi dXi + y 2 dX 2 + . . . + y. dX n (2.02)
where X\ 9 . . . X n are the geometrical coordinates and yi, . . . y n the
generalized forces. If the work is due in part to nonmechanical
forces (electric, magnetic, capillary, etc.), the form of this expression
is not changed : there only appear additional terms of the same type.
Of particular importance to us is the form which the element of
work takes in simple systems (section 2). The only force with which
such a system can act upon the outer world is the pressure upon its
boundaries. The pressure is the force per unit surface of the wall and
has the direction of the normal to it. Consequently, the force acting
upon an element d2 (Fig. 3) of the surface is F = pd2 in the direction
n. Let us suppose that, because of the pressure in the system, its
surface gets slightly displaced from the position AB into A'B'
so that the element dS moves into the new position d2' through
the very small distance A/. The work of the force upon this
element is then, according to the expression (2.01), equal to
dSA/ cos(n,A/), and the work of the forces of pressure upon all
the surface to DW = p f dSA/ cos(n,A/). In the limit, when A/ is
infinitesimal, every element, in being displaced from dl, todS', sweeps
18
117
THE FIRST LAW OF THERMODYNAMICS
19
over an oblique cylinder whose volume is, precisely, dA/ cos (n,A/).
The integral f dSA/ cos(n,A/) is, therefore, the total increase of the
volume of the system dV:
(2.03)
We shall, in general, denote by capital letters quantities referred to
the whole system (N mols) and by small letters quantities referred to 1 mol
of it. w represents, therefore, the molal work (W = Nw, N being the
mol number), or
Dw  pdv. (2.04)
It is work done by the system : We define the sign in such a way that
work of this kind is counted as positive, while work done by outer
FIG. 3. Element of work
in simple system.
FIG. 4. Graphical repre
sentation of work.
o v
FIG. 5. Work in cyclic
process.
forces against the system is considered as negative. The expression
(2.04) receives a simple graphical representation in the (p, v) diagram
(Fig. 4). Let the points A, A' correspond respectively to the initial
state of the system and to its state after an infinitesimal expansion dv.
The product pdv = Dw is, then, represented by the infinitesimal area
under the element AA' y the shaded strip of our diagram. If a finite
process of expansion is such that it can be represented by a curve
of the (p, fl)diagram, for instance, by the segment AB, the total
work (per mol) done in it by the system is equal to the integral of the
expression (2.04)
w
f
* A
pdv.
(2.05)
Its graphical representation is, obviously, the total area under the
segment AB. Since the work w is, in this case, positive, Fig. 4 permits
us to formulate the following rule. An area represents positive work
when it lies to the right of the curve giving the process (for the observer
looking in the direction of the process, i.e. from A to B). This rule is
20 TEXTBOOK OF THERMODYNAMICS II 8
borne out also in the case of a cyclic process where the system returns
to its initial state (Fig. 5). While expanding (ACB), the system does
positive work; while contracting (BDA), negative. The negative
area (double shaded) is to the left of the curve BDA and must be sub
tracted from the positive area under A CB. The difference is the net
(positive) work done by the system; it is represented by the area
enclosed by the circuit ACBD and lying to the right of it.
Exercise 8. (If the force is measured in dyne /cm 2 and the volume in cm 3 , the
work is expressed in erg). Calculate the work of vaporization of water. When
1 mol of water (18 grams) is vaporized at 100 C, the increase of volume is VB VA =
30186 cm 8 mol" 1 . The pressure in the process is constant and equal to 1 atm:
p = 1.013 X 10 6 dyne/cm 2 .
Exercise 9. Calculate the work necessary to compress 1 mol of a gas from the
normal state (VA = v = 22414cm 8 , PA = 1 atm) to half the normal volume
(VB = fo), assuming that the process follows the law of BoyleMariotte: pv =
8. Linear differential expressions and exact differentials. It will
be well to recall to the reader a few mathematical facts which, although
simple, are very fundamental in thermodynamics since the whole of
this science may be said to be built up on them. Mathematically
speaking, formulas (2.01) to (2.04) are examples of linear differential
expressions, whose general form is
k(xi ' Xn)dXl "

For instance, in the particular case of only two variables x,y
it reduces to
L(x,y) = M(x,y)dx + N( Xj y)dy. (2.07)
The most interesting question which arises with respect to these
expressions is whether they are " exact differentials 1 ', meaning whether
they can be obtained by differentiating some function f(xi, . . . x n ) of
the same variables. According to the rules of differentiation the total
differential of / is
dx k . (2.08)
We see, therefore, that the linear differential expression (2.06) is
an exact differential when
Jf*j, (*l,2...n); (2.09)
9**
II 8
THE FIRST LAW OF THERMODYNAMICS
21
in words, when the coefficients Mk are the partial derivatives of a
function/(#i, . . . x n ). If we single out the two independent variables
Xk, xi the second partial does not depend on the order of differentiation
Fropl'the UJfmulas (2.09) there follows, therefore,
; , (ft,/  1, 2 ...),
(2.10)
the socalled reciprocity relations. It is clear that the validity of the
reciprocity relations is a necessary condition for the expression (2.06)
being an exact differential. Moreover, it is shown in calculus that
this condition is also sufficient: Whenever the relations (2.10) are
satisfied, a function can be found such that the coefficients Mk are its
partial derivatives. This function is called
the potential of the expression.
We conclude from this that a linear
differential expression of the type (2.06) is
not, in general, an exact differential. If the
functions Mk(xi, . . . x n ) are selected at ran
dom, they will, in general, not satisfy the
reciprocity relations. The exact differential
is, therefore, a rare exception.
Of great importance is the concept of the p IG>
line integral of a linear differential expression.
To fix our ideas, we shall explain this notion
for the case of expression (2.07) with only two variables x, y, which we
shall interpret as the cartesian coordinates of a plane (Fig. 6). The
line integral along the curve from A to B
[M(x,y)dx + N( X ,y)dy}
is constructed in the following way. The first element is obtained by
substituting, for the arguments of the functions M and N, the values
of x and y in the point A and, for dx and dy, the increments which lead
from A to an infinitely close point A 1 on the curve. For the second
element one takes the values of M and N at the point 4' and the
increments leading to the point 4", and so one moves from point to
point along the curve, until its end B is reached. The integral is the
sum of all the elements so constructed. In the case of n variables they
6. Notion of line
integral.
22 TEXTBOOK OF THERMODYNAMICS II 8
must be interpreted as the coordinates of an wdimensional hyper
space, and the generalization is obvious.
When the linear differential expression is an exact differential (2.08),
its line integral can be given explicitly. For instance,
f
J A
The line integral depends, then, only on the coordinates of the initial
point #i A) , . . . x (A) and of the final x{ B \ . . . x ( * } and not on the path
between them. It has the same value for all the curves going from
the same initial to the same final point. In particular, when the curve
is closed, the initial and the final points coincide and the expression
(2.11) is equal to zero:
= 0. (2.11)
If V^ Mkdxk is an exact differential, its line integral over any closed
11
circuit vanishes. It is shown in mathematics that the inverse theorem
is also true : If the line integral of a linear differential expression vanisjte$
for every arbitrarily chosen closed circuit, it is an exact differential.
To summarize, we can say that the following three possible prop
erties of the linear differential expression (2.06) are completely equiva
lent: (1) It is an exact differential; (2) its coefficients satisfy the
reciprocity relations (2.10) ; (3) its line integral vanishes for any closed
curve. If the expression has one of these three properties, it has also
the other two.
Supposing that the linear differential expression (2.06) is not an
exact differential, the further question arises whether it can be trans
formed into one by multiplying it by a suitable function X(*i, . . . x n ).
If this is the case, we say that it admits an integrating multiplier or,
simply, that it is integrable. After multiplication by X the coefficients
of the expression become XM*, and, if it is integrable, there must be
satisfied, instead of (2.10), the reciprocity relations
In the case of the differential expression (2.07) with two variables,
there is only one reciprocity relation, 3(XAf)/3y = 9(XW)/9#, which
can be regarded as a partial differential equation with X as the depend
ent variable. Since such an equation has always a solution, a linear
H9 THE FIRST LAW OF THERMODYNAMICS 23
differential expression with two variables is always integrable. This is
no longer the case if n is larger than 2. For instance, when n = 3,
there exist the three relations
9*3 9*2 ' 9*i 9*3 9*2 9*i
of which two are independent. In general, it is impossible to satisfy
two equations by one function. This is brought out by the fact that
X can be eliminated from the three relations which leave as the result
of this elimination
,, ,_ /
Mi [  ]+M 2 [  f+Mz I ~  1 = 0. (2.13)
9*2 / \9*i 9*3 / \9*2 9*i
(The simplest way of checking this is to substitute the form of the
coefficients M k = (9//9**)A following from (2.12). This equation is
known as the condition of integr ability. Unless it is satisfied the dif
ferential expression does not admit an integrating multiplier. In the
general case of n variables, there must be satisfied a condition of the
type (2.13) for each triple of variables.
It should be noted that the integrating multiplier, when it exists,
is not unique. In fact, suppose that the linear differential expression
(2.06) is transformed by multiplication by X into the exact differential
df(xi 9 . . . x n ). The further multiplication by any function of/, say
<p(f), changes it into <p(f)df, which is also an exact differential. There
fore, X^>(/) is another integrating multiplier.
Exercise 10. Check by means of eq. (2.12) that L(p, v) = pdv vdp admits
an integrating multiplier, namely X = 1/t; 2 .
Exercise 11. Decide by means of the criteria (2.10) and (2.13) whether the
following linear differential expressions are exact differentials, integrable, or non
integrable:
(a) (2xydx + x*dy
(b) zdx + xdy +]ydz,
(c) yzdx + zxdy f xydz,
(d) (y + z)dx + (z+ x)dy + dz
9. Application to the element of work. In order to apply these
mathematical theorems to the physical concepts of section 7, we have
to find out whether in a simple system the element of work, Dw = pdv,
is an exact differential. As pointed out in section 3, the state of a
simple system is completely determined by two of the three variables
24 TEXTBOOK OF THERMODYNAMICS II 10
p, v, T. Let us use for its description T and v: The fact that dT does
not appear in the linear differential expression Dw means, then, that
the coefficient of dT is zero. The expression should be written
Dw SB pdv + Q.dT, and it is easy to see that it does not satisfy the
reciprocity relation (2.10) or dp/$T = 30/90, since the left side is
equal to the coefficient /3i (1.04) and is, generally, finite and the right
is 0. We conclude, therefore, that the element of work is not an exact
differential. This is confirmed by considering the line integral of Dw
over a closed circuit: we found in section 7 that it is represented by
the enclosed area while, in the case of an exact differential, it would
be equal to nothing. The same applies to the general case of a system
with n degrees of freedom. The difference between the element of
work (2.02) in mechanics and in thermodynamics is that, in mechanics,
the generalized forces y k are supposed to depend only on the coordi
nates of position Xi, . . . X n , whereas, in thermodynamics, they are,
in addition, functions of the temperature T.
y k = y k (Xi, X 2 ,... X n , T), k = 1, 2, . . . n. (2.14)
The term with the differential dT has the coefficient zero, and,
therefore, it is impossible to satisfy some of the reciprocity relations,
just as in the case of the simple system. This is the reason why we
use for the element of work the symbol Dw: we wish to reserve the
symbol d for exact differentials.
A system can be considered as defined only then when the forces it
exerts on its environment are known. In other words, eqs. (2.14)
must be considered as given. As they are completely analogous to
the equation of state, p = p(v,T), of the simple system, we conclude
that a homogeneous system has one less equation of state than it has
independent variables. (Compare also section 27).
10. Heat and heat capacity. As the invention of the thermometer
led to the concept of temperature (section 1), so another physical
instrument, the calorimeter, helped to develop the important concept
of heat. We have already seen that, when two systems of different
temperatures are brought in contact, the warmer cools down and the
colder heats up until equilibrium is reached. A calorimeter is a
standard system with which different bodies are brought in contact
in order to study their respective losses of temperature in relation to
its gain. To fix our ideas, let us suppose that the calorimeter consists
of one kilogram of water (initial temperature T\) into which the bodies
experimented upon (initial temperature Ti > 7i) can be dropped,
after which the combined system assumes a final temperature T inter
II 10 THE FIRST LAW OF THERMODYNAMICS 25
mediate between the two (T% > T> 7\). The notion of heat was
evolved from such measurements in the eighteenth century when the
calorimeter technique was not yet very precise. Within this limited
accuracy the following result was obtained: In experiments with
the same body, the final temperature T adjusts itself for all variations
of the initial temperatures T\ and T% in such a way that the ratio
rip rr*
C = 2 _  constant. (2.15)
T TI
C is, therefore, a characteristic constant of the particular body or
system and is called its heat capacity. Comparing different bodies of
the same material, it was, further, found that their heat capacities are
proportional to their masses. For a body of the mass Af, one can
write C = Me', where c f is the heat capacity per unit mass or the
specific heat of the material. Pouring into the calorimeter an additional
kilogram of water, we can determine the heat capacity C w of this
water like that of any other system. In the units implied in
formula (2.15) it turns out to be 1, and this is also the heat capacity
of the calorimeter itself, which was taken to consist of the same
amount of water (C w = 1). Equation (2.15) can, therefore, be
rewritten as follows
C (T  r 2 ) + C w (T  Ti) = 0. (2.16)
This relation has the form of an equation of conservation: It sug
gests that there is something that does not change in the process of
calorimetric conduction. We call this invariable function or quantity
heat. In fact, both terms of the left side have the form
Q = C(T  T), (2.17)
where T' is the initial and T the final temperature. The first term
relates to the body experimented upon and is negative; it represents
the heat lost by that body. The second, oppositely equal to the first,
is the heat imparted to the calorimeter.
The modern improved measurements have completely borne out
the fact that there is conservation of heat in all processes of conduction.
However, its definition in terms of the temperature had to be refined
somewhat, since it was found that the heat capacity is not strictly
constant but depends on the temperature interval. If Q is the heat
imparted to a system when its temperature is raised from T' to T,
we define
C Q/(T  n (218)
26 TEXTBOOK OF THERMODYNAMICS II 10
as the mean heat capacity for this interval. When the interval is
small, T = T  Ar, this can be written C = A^/Ar, and in the
limiting case of Ar = it becomes
in
(2.19)
the heat capacity at the temperature T. (We prefer the symbol D be
cause DQ is not an exact differential, as we shall see in section 13).
Multiplying this by dT and integrating, we find the more accurate
expression for the heat imparted to the system
/
CdT, (2.20)
which takes the place of the approximation (2.17) and must be, also,
substituted into the equation of conservation (2.16):
L
T s*T
CdT+ I C w dT = 0. (2.21)
J Tl
The accepted unit of heat is the (gram} calorie, the heat which
must be imparted to 1 g of water in order to raise its temperature by
1C. There are several different definitions of it in use, the most
important being: (a) The 15 calorie, the heat necessary to warm 1 g
of water from 14.5 to 15.5 C. (b) The mean calorie, one one
hundredth of the heat which raises the temperature of the same mass
of water from to 100 C. The latter is larger than the former in
the ratio 1.00024. One thousand gramcalories form one kilogram
calorie.
As to specific heats, we shall refer them, in the case of chemically
pure substances, to 1 mol of matter and denote by c the molal heat, i.e.
the heat which raises the temperature of 1 mol of the material by 1 C.
This definition is incomplete because the specific heat depends also
on the pressure conditions during the measurement; we shall have to
say more about this in section 14. Some authors use two terms: atomic
heat (heat capacity of 1 mol of a monatomic substance) and molecular
heat (1 mol of a chemical compound), but we shall apply the single
term, molal heat, to both cases.
Joseph Black, who established the " doctrine of specific heats,"
discovered also another phenomenon, the latent heat of transformations.
He observed (1761) that during the process of melting of ice (while
the temperature of the waterice mixture remains constant) a definite
amount of heat must be imparted to it, to transform it from the solid
II 11 THE FIRST LAW OF THERMODYNAMICS 27
state into the liquid, which he called the latent heat of fusion. Later
he also proved the existence of a latent heat of vaporization and approx
imately measured it. We shall denote the total latent heat by L and
the molal latent heat by /.
Chemical reactions are also accompanied by a positive or negative
heat development called the heat of reaction. If heat must be imparted
to the system in order to make the reaction proceed, it is called endo
thermic; in the opposite case, exothermic. We shall adhere to the con
vention that Q is positive when it is imparted to the system. Therefore,
the heat of an exothermic reaction will be counted as negative, of an
endothermic as positive.
In modern usage, the phenomena of change of volume and pressure
with temperature (treated in Chapter I) are spoken of as the thermal
properties of matter, while heat capacities, latent heats, etc., belong to
the caloric properties of matter.
Exercise 12. Into a calorimeter consisting of 200 g of water (initial temperature
/i =s 20 C) are thrown pieces of metal weighing 300 g and having the initial tempera
ture /2 = 100 C. The temperature of equilibrium / is for (a) aluminum (ju = 27.0)
39.5C, (b) iron (/z = 55.8) 31.9C, (c) zinc (ju  65.4) 29.8 C, (d) silver ( 
107.9) 26.2 C. Find the specific heats per 1 g and the molal heats of these four
metals, (/z denotes the atomic weight.)
Exercise 13. The molal latent heat of fusion of water (/z 18) is / 1440 cal
mol" 1 . We drop into a mixture of 1000 g ice and 1000 g water 300 g of copper
(/u = 63.6, molal heat c = 5.73 cal mol"" 1 ). How much ice will melt before the
system comes to equilibrium at C?
11. History of the first law. The notion of energy and of its con
servation was, first, established as a theorem restricted to the science
of mechanics. Leibnitz showed in 1693 that, in an isolated mechanical
system, there remains constant the sum of the kinetic and the potential
energies, called by him the live and the dead forces (vis viva and vis
mortua). If the system is not isolated, these energies are convertible
into work done against outer forces. About seventy years later was
formulated the law of the invariability of heat in all processes of con
duction, as stated in the preceding section. The first law of thermo
dynamics asserts that mechanical energy and heat, the two quantities
subject to laws of conservation, are not generically different, but
equivalent to each other and interconvertible.
A vague suspicion that mechanical action can have heating effects
seems to have existed since early times, but the first precise statement
that mechanical work can be converted into heat and the first attempts
to put this theory upon an experimental foundation were due to
Benjamin Thompson, a native of Connecticut, better known under
28 TEXTBOOK OF THERMODYNAMICS II 11
his Bavarian title of Count Rumford (1798). Supervising the boring
of cannons in the arsenal of Munich, Rumford was struck by the high
temperature which the brass shavings attained and by the general
large development of heat in the process. He could think only of two
causes, other than the conversion of work into heat, which, conceivably,
could have produced this effect: On the one hand, there may exist,
he thought, a latent heat of breaking up solid brass into chips,; on the
other, the borer may induce some chemical action of the air upon
brass attended by heat development. The first possibility was dis
proved by special measurements of two different kinds. One should
expect that the surmised latent heat is due, largely, to a difference in
the heat capacities of the solid metal and of the shavings. However,
no such difference could be detected by calorimetric determinations
especially undertaken for the purpose. Still more conclusive was a
second line of reasoning: When a blunt borer rotating with much
friction was substituted for the sharp one, the production of chips
decreased many times while the yield of heat remained about the
same. As to the possible action of air, Rumford tried to meet this
objection by immersing the whole apparatus in water so that the air
had no access to any of the heated parts. This change of conditions
had no appreciable effect on the rate of the heat development.
Rumford gives very explicit data how much heat the power of a live
horse produces per hour. Translated into modern units, they are
estimated to mean that 1 (gram)calorie is equivalent to 5.45 X 10 7
erg, a result which is 31% above the value at present accepted. A
sequel to Rumford's work was a paper by his associate in the Royal
Institution of London, Sir Humphry Davy (1799). In order to refute
the objection that the heat, in the above experiments, could have
come from the water (which might have exercised upon brass a chem
ical influence similar to that of air), Davy studied the conversion of
work into heat in vacuo. By means of a clockwork he rubbed two
pieces of ice against each other under the jar of an air pump. The
effect was unmistakable, inasmuch as the ice melted faster with
rubbing than without, but no quantitative determinations were
attempted.
The brilliant lead of Rumford and Davy was not followed up by
other investigators. Forty years passed as a period of quiescence dur
ing which these problems were, practically, never mentioned as far as
the literature of physics and chemistry is concerned. Not until the
decade from 1840 to 1850 was the principle of conservation of energy
established and accepted, owing to the independent work of several
men belonging to different nations. The French engineer S6guin
II 11 THE FIRST LAW OF THERMODYNAMICS 29
(1839) formulated the theory of the equivalence of work and heat with
great clarity and carried out much experimental work trying to prove
it. The idea of these experiments was excellent and amounted to a
reinterpretation of the formula of adiabatic expansion from the point
of view of conservation of energy (compare section 18). Their execu
tion and analysis, however, were imperfect, and Seguin himself
admitted that his results succeeded only in making his contention
probable and did not constitute a convincing proof. Quite independ
ently and with equal lucidity the principle was stated by the German
physician Robert Julius Mayer (1842), who also computed the ratio
between the units of heat and work (mechanical equivalent of heat)
from data on the specific heat of gases then available (compare section
15). This method, although not quite free from objection, disclosed
an unusual insight into the nature of the phenomena involved in it.
The result at which he arrived is 1 calorie = 3.57 X 10 7 ergs, or 15%
below the correct value. However, Mayer seems to have overlooked
that even the most ingenious hypothesis needs an experimental
demonstration in order to become a scientific law: no experiments
were either planned or suggested by him. The credit of having set
the principle of conservation of energy upon a firm experimental
foundation is, entirely, due to James Prescott Joule of Manchester
(England). Owing to delicate health, Joule did not receive a regular
course of instruction but was tutored at home; among his teachers
was the chemist, John Dalton. Joule conceived the idea of the
equivalence of heat and work at an early date and began his experi
ments upon it in 1840 at the age of 20. They were carried through
in his private laboratory built for him by his father, a retired brewer.
He was led to the problems of conservation of energy from a study of
electric currents, having ascertained (in 1840) that there is a constant
relation between the work necessary to maintain a current and the
heat generated by it (Joule's heat). In his first determination of the
mechanical equivalent of heat (1843) he utilized just this fact: The
current was supplied by an electromagnetic generator and passed
through a resistance immersed in a calorimeter. While this circuit
operated at a stationary rate, he determined, on one hand, the work put
into the generator, on the other, the heat developed in the calorimeter
(result: 1 cal = 4.6 X 10 7 erg). Best known is his direct determina
tion of 1849 in which he produced heat by churning water and other
liquids with paddlewheels, and which gave 1 cal = 4.154 X 10 7 erg
(for a long time, the accepted standard value, unsurpassed for several
decades), but his work included the study of a great variety of processes
relating to the interconvertibility of mechanical, thermal, electrical,
30 TEXTBOOK OF THERMODYNAMICS II 11
and chemical energies. As one of the founders of the principle of
conservation of energy must be also regarded Hermann Helmholtz,
at that time a young surgeon in the Prussian army. He at once recog
nized its importance and scope and was instrumental in bringing it
to the attention and acceptance of scientists by publishing a brilliant
paper (1848) in which he followed its workings through numerous
applications to the sciences of physics, chemistry, and physiology. 1
It is very remarkable that none of these four men was by training
a physicist. Sguin was brought up as an engineer, Mayer and
Helmholtz as physicians, while Joule having received an irregular edu
cation cannot be classified with any professional group. The significant
fact that the contemporary physicists had no part in the development
of a most important phase of their science cannot be denied, even if
Helmholtz and Joule later became ornaments of their profession.
The reason for this has been pointed out by Ostwald and others: It
was the time of reaction after the collapse of Schelling's " Natur
philosophie". All the physicists had been, for a longer or shorter
period, under the influence of that movement. Now they were tired
of fruitless generalities and anxious to make amends for the wasted
time by conscientious factual research. They became suspicious of
anything that was reminiscent of a speculative anticipation of nature.
This frame of mind accounts for the hostility with which the mechan
ical theory of heat was received in physical circles. On the other hand,
the active interest developed by engineers on its behalf hardly needs
any explanation : The production of power from fuel is the engineer's
everyday pursuit which compels him to think continuously about
the connection between heat and work.
These facts help us to understand the aloofness of the physicists
and the active participation of an engineer in investigating the equiva
lence of mechanical and caloric energy. It remains, however, unex
plained how this principle could enlist the interest of two distinguished
medical men. The writer began to feel that there might lie a historical
problem in this and that it would be worth his while to inquire into the
interests and the scientific background of physicians and physiologists
of the period between 1820 and 1840. These suspicions were confirmed
1 Posthumous notes of S. Cannot disclose that this author, who wrote his " Reflex
ions sur la puissance motrice du feu " strictly from the point of view of the indestruc
tible caloric fluid, in later years (before his death in 1832) became more friendly to
the idea of the equivalence of heat and work. However, it is hardly possible to
assign him a share in the discovery of the first law, on the strength of this, because
he did not possess either a rigorous formulation of it or any new empirical evidence
to support it. Besides, his notes were published for the first time in 1878, a quarter
of a century after the law had ceased to be controversial.
II 11 THE FIRST LAW OF THERMODYNAMICS 31
by the following passage in one of Joule's papers: " On conversing a
few days ago with my friend Mr. John Davies, he told me that he had
himself, a few years ago, attempted to account for that part of animal
heat which Crawford's theory had left unexplained, by the friction
of the blood in the veins and arteries, but that finding a similar
hypothesis in Haller's 'Physiology 1 , he had not pursued the subject
farther " (Postscript to the paper in Phil. Mag. 23, 1843). These
lines tell us about a third physician who was familiar with the idea of
the equivalence of heat and work. Moreover, they give us a hint
that the source of his interest in the subject lay in the theory of animal
heat, with a reference to Haller's physiology as a clue.
At present, the name of Albert Haller may not be generally
known, but in the eighteenth century he was considered a scientific
colossus and enjoyed an international prestige, as an authority in
medicine and botany, which is equaled only by that of Newton, in
his sphere and at his time. A native of Berne (Switzerland), he became
the main scientific figure at the university of Gottingen. In German
speaking countries his reputation was further enhanced by important
poetical writings. His " Outline of Physiology " appeared in 1747 and
was immediately translated into French and English. It is not too
much to say that it marked a new epoch in that branch of science
and became the main influence in the education of several generations
of physicians. The last German edition was published in 1822, and
we have seen that Dr. Davies consulted this treatise even about the
year 1840. The passage which Dr. Davies must have had in mind is
contained in section 303 of the first volume. 1 On the preceding pages
Haller analyzes the uses of the lungs and comes to the conclusion that
the animal heat is produced in this organ and imparted to the blood
passing through it. Of considerable interest to us is his hypothesis as
to how the heat originates: In his own words, it arises " from the
alternate extension and contraction, relaxation, and compression of
the pulmonary vessels, by which the solid parts of the blood are per
petually rubbed together and closely compressed in the attrition that
is made during expiration, as it is more rapidly moved and ground
together during expiration". Very important is, also, the following
sentence: " Nor is it any objection to this that water cannot be made
hot by any friction. Nor in reality is this assertion true; for water by
violent winds and motion, as well as milk, acquires some degree of
warmth". These quotations contain Haller's theory of respiration
which, for a long time, dominated the science of medicine. It is quite
a surprise to find that this theory was essentially based on the idea
1 Not section 304 of Vol. II, as Joule gives the reference.
32 TEXTBOOK OF THERMODYNAMICS II 11
that heat can be generated by mechanical work, although it was
enunciated 50 years prior to the work of Rumford. 1
We have found a convincing proof of the fact that, at a certain
epoch, the physiologists were much concerned with the problem of
the relation between work and heat, and we shall see below that they
were keenly aware of the work of Rumford and Davies, seeing in it a
confirmation of Mailer's hypothesis. It is surprising, at first sight,
that this interest could have persisted until the middle of the nineteenth
century, since the modern combustion theory of respiration, suggested
already by Priestley, received strong support from the measurements
of Lavoisier and Laplace (about 1781) and was rediscovered indepen
dently by the English physician Crawford (referred to by Joule as
Crawford's theory). A closer investigation shows, however, that there
were excellent reasons which prevented the acceptance of the combus
tion theory as a complete explanation of animal heat. Lavoisier and
Laplace had found that the oxidation of carbon contained in the car
bohydrates of the blood is insufficient to account for the developed
heat, and they expressed the conjecture that the balance is produced
by the oxidation of hydrogen. An experimental test of this hypothesis
required measurements extremely delicate and difficult for the prim
itive technique of those times. A long time passed, therefore, before
they were even attempted. At length (in 1821), the Academic des
Sciences announced a prize for the investigation of this problem, and
this induced Dulong and Despretz to start work on it. Despretz 's
essay was submitted to the academy in due time and received the prize
on June 1, 1823 (published in the Ann. Chem. Phys. (2) 26, p. 337,
1824). He concluded from measurements on 200 animals that the
complete heat of oxidation of both carbon and hydrogen accounts
for the larger part of the animal heat but not for the whole of it.
There remained a defect of about 20% for the explanation of which
Despretz fell back upon Haller's theory, saying that it could be
produced " by the motion of the blood and the friction of its different
parts". The results obtained by Dulong were quite in agreement with
those of Despretz's. He read his paper before the academy in 1822,
but he did not contend with it for the prize and he resisted the urging
of his friends to have it published (appeared after his death: Ann.
Chim. Phys. (3) 1, p. 440, 1841). The reasons for this strange reluc
tance became clear when his posthumous papers were examined:
Dulong had no faith in the thermochemical data on which his own
and Despretz's conclusions were based. He intended to redetermine
1 Even Haller cannot claim priority in this matter, since similar views were held
in the seventeenth century by the latroPhysical School.
II 11 THE FIRST LAW OF THERMODYNAMICS 33
the heats of formation of carbon dioxide and of water in order to
revise his essay. This rede termination, finally carried out, was the
last scientific work done by Dulong: he obtained new and more
accurate data but died before bringing his results into shape for
publication (posthumously edited by Cobart, Ann. Chim. Phys. (3) 8,
p. 183, 1843). While his work confirmed the previously accepted
heat of formation of carbon dioxide as fairly accurate, it established
an entirely new value for the heat of formation of water, showing
that the old data were much too low. This correction was more than
sufficient to remove the discrepancy between the animal heat and the
chemical heat development of respiration and constituted a conclusive
proof of the oxidation theory. However, it must be remembered that
this knowledge came rather late: between 1823 and 1843 the last
word of physiological science was the existence of a defect of animal
heat which, apparently, was left unaccounted for by the combustion
theory.
There remains the question how widely this knowledge was diffused
among physiologists and physicians. The writer happened to come
across a very comprehensive German " Handbuch " of physiology of
that time written in six volumes by leading authorities under the
editorship of Karl Friedrich Burdach. The account of the theory of
respiration, due to Ernst Burdach, is contained in the sixth volume,
which appeared in the year 1840. Many pages are devoted to the
history of the two theories, and both sides of the question are discussed
at considerable length. After a lucid presentation of Haller's views
(p. 544), Burdach turns to " the doctrine of the resemblance of respira
tion and combustion". It appears from his account that in the period
between 1820 and 1840 the physicians interested in the theory of
respiration were preoccupied, like Joule's friend Dr. Davies, by
" that part of animal heat which Crawford's theory left unexplained".
Most of them pointed to Haller's theory as to the most likely explana
tion and maintained that the origin of the unaccounted heat lay in
the mechanical action of friction in the arteries. These mechanical
effects were, in their turn, supposed to be produced by the " vital
principle " (puissance vitale) invented by Barthez. Burdach gives
many references to authors who advocated this view. Among them
were some of the leading men of science, like Despretz and Treviranus,
but also some of the rank and file. For instance, the writer had the
opportunity of reading in the originals a German pamphlet by F. Lau
(Bonn 1830) and an English paper by J. M. Winn (Phil. Mag. 14,
p. 174, 1839). They are of great interest as an indication into how
wide a circle the problem had spread. Lau and Winn were, apparently,
34 TEXTBOOK OF THERMODYNAMICS II 11
ordinary practitioners without much knowledge of the literature of
the question or understanding of its deeper implications. Yet, their
theories were only modifications of Haller's views. It is interesting
to note that Burdach himself, who was not in sympathy with these
explanations, did not question the fact that heat can be produced by
friction, but regarded it as a matter of common knowledge, citing, in
this respect, Rumford and Davy. He only doubted whether this
cause would be sufficient to account for the desired effect quantitatively.
The above facts throw a new light on the role of the medical
profession in the history of the principle of conservation of energy.
There was, actually, no break of tradition between the work of the
eighteenth and the nineteenth centuries. However, this tradition
was handed down through the unexpected channel of the science of
physiology, whose representatives acted, for half a century, as the
custodians of the idea of equivalence of heat and mechanical work.
If this statement implies that the historical role of Mayer and
Helmholtz consisted in turning back this principle to physics, the
branch of science where it naturally belongs, this is in no way a
detraction from the merits of these two great men: They were not
merely passive transmitters of a concept, because there is an enormous
gap between the vague knowledge current among physicians, that heat
is generated by mechanical action, and the scientific formulation of
the principle of equivalence, which could be bridged only by men of
unusual ability. Helmholtz, eventually, developed into one of the
greatest scientists of all times, and Mayer showed in his later physio
logical writings (1845) a clearer understanding of the nature of animal
metabolism than any of his contemporaries. Nevertheless Mayer was
entirely unsuccessful in his endeavors to persuade the physicists and
chemists of the validity of his conception. The reason for this was,
without doubt, that he could not offer them much in the way of an
experimental proof at a period when they were particularly reluctant
to accept anything on insufficient empirical evidence. A few years
latter Helmholtz had far better success because, in the meantime,
Joule's work had put the principle of conservation of energy on a
secure experimental foundation.
CHAPTER III
THE FIRST LAW OF THERMODYNAMICS AND THE
CALORIC PROPERTIES OF MATTER
12. Reversible and irreversible processes. It was pointed out in
section 2 that the thermodynamical description of a system, by such
variables as its temperature and pressure, implies that it is in equi
librium. In general, this equilibrium is disturbed if the system is
subjected to any physical process. For instance, if we exert a mechan
ical action upon it or bring it under the influence of a source of heat,
the result is a complicated transient state, of longer or shorter dura
tion, during which all the physical characteristics of the system are
different in every point of it. The quantitative description of such an
occurrence is, therefore, entirely outside the scope of thermodynamics.
It is possible to follow through by thermodynamical methods only
processes which take place so slowly that they can be regarded as a
succession of states of equilibrium. Strictly speaking, the change of
the system in a finite time must be even infinitesimal and the rate of
change infinitely slow.
Such infinitely slow processes have an important theoretical
advantage in that they are reversible, that is, they can be gone through
in both directions. In fact, let us take the example of a simple system:
If we change two of the three variables, say v and T, so slowly that
the system can be regarded as in equilibrium at any moment, the
equation of state applies all the time and the third variable p is also
completely determined. When we go through the same values of
v and T in the opposite direction, the succession of states of the system
is, simply, reversed. For every infinitesimal step, the work done by
the system is expressed by Dw = pdv. When the processis reversed,
dv changes its sign while p has the same values so that the work retains
the same absolute value as in the direct process,, but has the opposite
sign, i.e. is done by outer forces against those of the system. We shall
see in the next section that the heat received by the system in every
infinitesimal part of the process is also a linear differential expression
so that the same conclusions hold with respect to it: If a certain
amount of heat is imparted to a system in an infinitely slow process,
35
36 TEXTBOOK OF THERMODYNAMICS III 12
the same amount is withdrawn from it when the process is reversed.
We have restricted ourselves to simple systems, only, to fix our ideas :
All our results apply also to the most general thermodynamical system
because, for every phase of it, there exist equations of state of the type
(2.14) and the elements of work and heat have always the form of
linear differential expressions.
When we turn to processes of a finite speed, we can no longer assert
that they are reversible. As an example, we consider a gas in a
vertical cylindrical vessel with a piston as its top. The reversible way
of compressing such a gas would be to increase the load on the piston
infinitely slowly by adding infinitesimal weights. When a finite
weight is added at a time, the compression goes at a finite rate, but
owing to the lag in the propagation of stresses, the compression and
the pressure immediately under the piston will be higher than in the
remainder of the gas and higher than in the reversible process. To
produce the same reduction of volume, the outer forces have, there
fore, to do more work than in the infinitely slow case : A part of the
work is consumed by elastic waves, currents, and other wasteful
processes in the gas, the socalled phenomena of internal friction. On
the other hand, if we expand the gas back to the original volume,
proceeding at a finite speed, the same lag will cause the pressure at
the piston to be lower than in an infinitely slow expansion. The gas
will do, therefore, less work than in the latter process. It follows
from this that the compression and expansion, at a rate which is not
infinitely slow, are irreversible processes since the work which must
be applied in the compression is larger than that received from the
system in the expansion.
The extreme case of rapid expansion is that when there is no outer
force 1 all, for instance, when the piston is suddenly removed. Such
an expansion into the void takes place without the system doing any
work. It is quite obviously irreversible, since it is necessary to apply
work of outer forces to compress the system back to the initial volume.
Another example of an irreversible process is the phenomenon of
thermal conduction of which we have made use in defining the con
cepts of temperature (section 2) and of heat (section 10) : If we bring
in contact two bodies having a finite difference of temperatures, heat
flows, spontaneously, from the warmer to the cooler, but never in the
opposite direction. All processes of nature proceed at a finite speed,
and experience tells us that they are all accompanied by effects of the
nature of friction or conduction which make them irreversible. The
engineer would call these effects " wasteful' ', since they do not produce
any work, and the irreversible process " less efficient " than the
Ill 13 CALORIC PROPERTIES OF MATTER 37
reversible. Strictly speaking, the reversible processes are an ideal
which exists only in our imagination. However, this is enough to
make its discussion very useful; besides, many real processes are con
ducted at so slow a rate that they can be regarded as reversible within
the limits of experimental error. Only such processes can be mathe
matically described in thermodynamics, or represented by means of
curves in diagrams. If the system is not in equilibrium no unique
values of temperature and pressure can be assigned to it.
13. Analytical formulation of the first law. Internal energy.
The purport of the first law has been made sufficiently clear in Chapter
II. Its gist is that heat is convertible into mechanical work and other
forms of energy, and vice versa, in the ratio of the mechanical equiva
lent of heat: 1 15gramcalorie = 4.185 X 10 7 erg. By means of
this conversion factor we shall always express heat and work in the
same units. What is, then, the fate of an infinitesimal amount of
heat DQ, imparted to a system by bringing it in contact with a suit
able heat reservoir? Since thermodynamics deals only with states
which are infinitely close to equilibrium we need not consider its con
version into kinetic energy: The heat DQ will, in part, cause the
system to do the work DW against outer forces; in part, it will be
used to raise the temperature and, perhaps, to overcome the resistance
of inner forces attending a change in volume and in other inner prop
erties. We denote the portion utilized for the latter purpose by DU
and obtain the equation :
= DU + DW. (3.61)
Wejmow that the element of work DW is not an exact differential,
but we do not yet know what the elements DQ and DU a^e, in a
mathematical sense, and it will be our next objective to find this out.
This is the reason why we use the symbol D: According to the con
vention made in section 8, the symbol d is reserved for expressions
which, we are sure, are exact differentials.
Let us now consider a reversible process in which the heat Q is
imparted to our system by bringing it, successively, in contact with
an infinity of heat reservoirs differing in temperature by infinitesimal
amounts. For every step of this process eq. (3.01) holds, and we have,
simply, to integrate it from the initial state of the system (1) to its
final state (2) : Q = f* DQ, and, according to section 7, W =/f DW.
We find, therefore,
DU = Q  W. (3.02)
f
J
38 TEXTBOOK OF THERMODYNAMICS III 13
Of particular interest is the case when the process is cyclic so that
at the end of it the system is, precisely, in the same state as in the
beginning. The system itself does not, then, contribute anything to
the energy balance. If we compare the conditions before and after
the cycle is carried through, the only changes are outside the system.
Namely, the heat Q is withdrawn from the reservoirs and the work W
done against outer forces. The law of conservation of energy leads,
therefore, to Q W 0, instead of (3.02), whence
DU=Q, (3.03)
for any reversible cyclic process. As we have seen in section 8,
eq. (2.11), it follows from the property (3.03) of the integral that the
expression DU is an exact differential. This means that there exists
a certain function U of the variables defining our system whose total
differential d U is identical with the differential expression D U. This
function U is known as the internal energy, Every function which is
completely determined by the variables of the system we shall call a
characteristic function of it. The internal energy is, therefore, one of
the " characteristic functions " of thermodynamical systems.
Instead of eq. (3.01) we have to write, therefore,
DQ = dU 4 BW. (3.04)
The element DQ being the sum of an exact and a nonexact dif
ferential is itself not an exact differential according to the criteria of
section 8. The integral (3.02) over a reversible process takes now
the form
Q  U 2  Ui + W. (3.05)
In words: the heat imparted to the system is used, partly, to
increase its internal energy from the initial value U\ to the final Z/2,
and partly to do outer work, Though derived for a reversible process,
this equation has a wider range of application. In fact, suppose that
in an initial state of equilibrium the system has the internal energy Ui.
We subject it, then, to an irreversible process and, after it is over,
wait until the system and its environment are again in equilibrium.
^Suppose it is now found that the internal energy has become C/2, that
outer sources have lost the heat Q, and that the work W has been
done against outer forces. The law of conservation of energy leads
again to the relation (3.05). In this sense it can be applied also to an
Ill 14 CALORIC PROPERTIES OF MATTER 39
irreversible process. In the particular case when the process is cyclic,
the initiaT and the final state are identical, J7 2 Ui = 0, and
Q  W. (3.06)
The gist of the first law of thermodynamics is that work can be
produced only at the expense of heat or some other form of energy.
Therefore, it can be formulated as the impossibility of a perpetual
motion machine (perpetuum mobile), as those delusive devices are
called which claim to do useful work, continually, without drawing
on any energy source. In its application to purely mechanical engines,
this principle had already been recognized by Stevinus, Newton,
and Leibnitz. The discoveries of the middle of the nineteenth century
(section 11) extended it to engines of any kind.
Exercise 14. 300 g of mercury are dropped from a height of 600 cm into a
calorimeter consisting of 500 g of water at 10 C. In this process the potential
energy of the mercury is, first, converted into kinetic energy and then into heat.
What will be the final temperature of the water and mercury? (The specific heat
of mercury is 0.0334 cal g" 1 deg" 1 , the acceleration of gravity g  980.7 cm sec" 2 ).
Exercise 15. The latent heat of vaporization of water is /  9730 cal mol~ l
(per 18 g). What part of it is spent on the work of expanding the steam, as calcu
lated in exercise 8?
14. Molal heat at constant volume and at constant pressure.
We are going to apply the first law, as formulated in eq. (3.05), to the
special case of a simple, chemically pure system. We shall refer
Q, U, W to one mol of the material of the system, denoting the molaj
quantities, by the small letters q, , w. Moreover, according to (2.04),
the element of work is, for a simple system, Dw = pdv, whence
Dq = du + pdv. (3.07)
The molal heat is, according to the definition (2.19) of specific
heats, given by
c  Dq/dT (3.08)
and depends upon the conditions under which the element of heat Dq
is imparted to 1 mol of the system. Two cases are of special prac
tical interest: (1) the volume of the system is kept constant (dv 0)
while heat is added to it; (2) its pressure is kept constant (dp ** 0)
The internal energy u depends on the variables which define the
system: in the case of a simple system, we can regard it as a function
of any two of the three parameters p, v 9 T. To treat the case of
40 TEXTBOOK OF THERMODYNAMICS III 14
constant volume it will be useful to describe it in terms of temperature
and molal volume:
u = u(T, v), (3.09)
a relation which is often called the caloric equation of state. Its total
differential has, then, the expression
which, substituted into (3.07), gives
If we put dv = 0, we obtain, comparing this with (3.08), the molal
heat at constant volume
c v =
On the other hand, to treat the case of constant pressure, we must
find an expression of Dq in terms of the differentials dT and dp. The
simplest way of doing this is to recall that, because of the equation of
state, v itself can be regarded as a function of T and p so that
Substituting this into (3.11)
Putting dp = 0, we find by comparison with (3.08) the molal heat
at constant pressure
The subtraction of (3.12) from (3.14) leads to the important relation
The reason why c p is larger than c 9 is obvious: If the system is per
mitted to expand while the element of heat Dq (per mol) is imparted
to it, this heat is used, not only to raise the temperature of the systen f
Ill 15 CALORIC PROPERTIES OF MATTER 41
but also to do work, namely, work against outer forces to the extent
p(dv/dT) p dT and against inner forces of the system itself in the
amount (3u/3) r  (dv/dT) p dT.
In the case of condensed systems (liquids and solids) the coefficient
of expansion (3^/3 T) p is so small that it can be neglected for most
purposes. Within this approximation it is unnecessary to distinguish
between molal heats at constant volume and at constant pressure
(c p c v = c). We postpone a detailed discussion of the molal heats,
as to their experimental values as functions of temperature, until
Chapter XVIII, but it will be necessary to say something about the
approximate laws for permanent gases and for solids in the next two
sections.
Exercise 16. The approximate value of Ow/dtOr for N 2 , at C and 1 atm, is
0.0047. Estimate (dv/dT) p from the data of Table 1, and say by what percentage
c p c v deviates from R.
IS. The internal energy of perfect gases. The question in what
measure the internal energy of gases depends on their volume can be
decided in an experimental way. A very suitable experiment for this
purpose is the expansion of a gas into the void mentioned in section 12.
Provided that the system is adiabatically isolated, neither heat is
imparted to it in this process nor does it do any external work, so that
the terms Q and W in eq. (3.05) of conservation of energy vanish, and
it is reduced to /2 U\ = or, referring it to 1 mol of the gas,
Au = U2 ui = 0. If the expansion is small, we can write with the
help of (3.10) and (3.12)
or
If the internal energy is independent of the volume, (du/dv)r = 0,
there follows Ar = 0: the expansion into the void is not attended
by any change of temperature. It is possible to test this point by
conducting the process while the system is in contact with a calorim
eter> instead of being adiabatically isolated. Joule carried out the
experiment (1845) in the following form: two large, closed copper
vessels were connected by a short copper pipe containing a stop cock.
While the stop cock was shut, one of them was filled with gas at high
pressure, the other evacuated and both were immersed into the same
water calorimeter. When the stop cock was opened, the gas rushed
into the empty vessel without doing work, but no change of temper
42 TEXTBOOK OF THERMODYNAMICS III 15
ature of the calorimeter could be discovered for any of the gases used.
(A somewhat similar experiment, tried by GayLussac as early as
1807, had a quite different purpose). It may be concluded from this
that the dependence of the internal energy of permanent gases upon
their volume is small, but the large heat capacity of the calorimeter
reduced the accuracy of the experiment too much to say more. To
increase the accuracy, it is desirable to remove the calorimeter alto
gether and to measure directly the temperature of the expanded gas.
However, this is impossible in the original form of the experiment:
after the stop cock is opened, the gas is in a turbulent state with no
uniformity of pressure or temperature. It comes to rest only after a
period of time long enough to vitiate the results by heat losses through
convection. A modification which avoids these difficulties was sug
gested by William Thomson (later Lord Kelvin) and carried through,
jointly, by Joule and Thomson in a
MI NI M 2 Nz series of investigations extending from
 l p i' T i! B ! P 2' T 2 ["> 1852 to 1862  The y maintained a sta
hi C A 2 tionary stream of gas through a tube
FIG. 7. JouleThomson A \A^ (Fig. 7), obstructed in the middle
process. by a cotton plug C. Owing to the friction
in the plug, there was no turbulence in
the flow and the gas was in a homogeneous state on either side of the
obstruction. Before passing it, its pressure, molal volume, temper
ature, and internal energy were pi, vi, Ti, u\, afterwards p2, 22, T%, #2.
The difference of temperatures T% T\ was measured by a thermo
electric couple. Heat exchange with the environment was prevented
by protecting the tube with heatinsulating material. As to the
kinetic energy of the gas, and the friction losses in the plug, they are
proportional to the square of the velocity of the flow, which was chosen
so low that they were entirely negligible, compared with the other
energy items. Let us set up the energy balance for a portion of the gas
which is contained (at a certain moment) between the crosssections
MI and Af2, for the time during which 1 mol of the gas flows through
the tube. At the end of this time it will occupy the new position
between NI and N2 and we can apply to this process the eq. (3.05)
of conservation of energy which reduces here to C/2 U\ + W = 0,
since no heat is imparted to the gas (Q = 0). As a result of the dis
placement, 1 mol of the gas (with the internal energy #2) appears on
the right side of the plug and 1 mol (of the energy ui) disappears on
the left: Uz U\ = u% u\. On the right side, work is done by
this part of the gas in pushing forward the adjacent portions through
the volume v% at the constant pressure pz\ on the left, work is done
Ill 15 CALORIC PROPERTIES OF MATTER 43
against it while it is displaced through the volume v\ (pressure pi):
W = p2V2 p\v\. Our equation assumes, therefore, the form
2 + p2V2 = U\ + piVl. (3.16)
We shall postpone the analysis of this equation and of the results
of Joule and Thomson until section 28. Suffice it to say here that,
generally, the partial (du/Qv)T turned out to be small, but not equal
to zero. Comparing the different permanent gases at room temper
ature, it decreases in the order in which the gases become more perfect
(section 6) and is almost negligible for helium. It is, therefore, logical
to define the perfect gas as characterized by an internal energy entirely
independent of its volume: (3w/3z>)r = 0, so that, from (3.12) and
(3.10),
du  c v dT. (3.17)
This implies, of course, that c v is independent of v, because of the
condition (2.10). As to its dependence upon the temperature, the
experiments with real gases show, again, that it is the smaller, the
nearer the gas approaches in its equation of state the behavior of
perfect gases. In the definition of the perfect gas c v is, therefore,
taken to be constant. Integrating eq. (3.17), there follows the expres
sion of the internal energy of perfect gases
u  c v T + u , (3.18)
where u (the socalled " zero point energy ") is an integration con
stant. This formula gives a good approximation to the internal
energy of real monatomic gases, although it is not safe to extrapolate
it to extremely low temperatures (section 105). In the case of diatomic
and polyatomic gases, the approximation is fair only for a limited
range of temperatures (on both sides of the room temperature): At
very low temperatures the molal heat decreases because the rotational
motions of the molecules require less energy; at very high temperatures
it increases because the molecular vibrations begin to come into play.
In view of this situation it is well to give the expression also for the
case when the gas is thermally perfect (i.e. obeys the equation pv = RT)
but calorically imperfect, to the extent that its specific heat, being in
dependent of v, is a function of T
c, dT + ,. (3.19)
In the special case c, = const, this becomes identical with (3.18).
44
TEXTBOOK OF THERMODYNAMICS
III 15
In either case, eqs. (dv/dT) p = R/p and (9w/9^)r = hold, whence
from (3.15):
c p  c v = R. (3.20)
This is the equation which J. R. Mayer used for the determination
of the mechanical equivalent of heat, taking, for c p and c v , data
expressed in calories (per degree and mol) and, for R, in mechanical
units (compare section 12). Theoretically, his method was not free
from objection, since, at the time, there was no foundation for the
hypothesis that in the expansion of gases no energy is needed to
overcome internal forces. In practice, however, the later work of
Joule and Thomson showed that the formula is satisfied fairly well:
Mayer would have obtained a pretty accurate value of the equivalent
if good data of c v and c p had been available. The modern value of
the gas constant in caloric units is R = 1.9864 cal deg mol" 1 .
In the case of monatomic gases the molal heats are very close to
c v = 3R/2, c p = 5R/2, so that the ratio is 7 = c p /c v = 5/3. The
data for the more important gases at 15 C are given in Table 4.
TABLE 4
MOLAL HEATS OF GASES
Gas
7 = C P /C V
c P /R
(c p  c v )/R
Gas
7 = C P /CV
Cp/R
(c p  c v )/R
A
.668
2.52
1.00
C0 2
1.304
4.41
1.03
Ne
.66
N 2 O
1.303
4.44
1.03
Xe
.66
NH 8
1.31
4.49
1.06
rig
.67
CH 2
1.31
4.27
1.01
H 2
.410
3.44
1.00
H 2 S
1.32
4.35
1.05
N 2
.404
3.50
1.00
SO 2
1.29
4.89
1.10
2
1.401
3.52
1.00
CN
1.26
5.36
1.09
CO
1.404
3.50
1.00
C 2 H 2
1.26
5.02
1.03
NO
1.400
3.52
1.00
C 2 H 4
1.255
5.08
1.03
HC1
1.41
3.56
1.00
C 2 H
1.22
5.84
1.05
cu
1.355
4.10
1.08
A more detailed account of specific heats of gases will be found
in Chapter XVIII.
Exercise 17. For water vapor in the range between t = C and * = 650 C
Nernst gives the empirical formula
c p  8.62 + 0,002 / + 7.2 X 10 / 2 cal mol" 1 deg 1 .
Assuming that eq. (3.20) is satisfied, find c v and determine the increase of internal
energy c*o ~ o.
Ill 16
CALORIC PROPERTIES OF MATTER
45
Exercise 18. For CO 8 between 75 C and +20 C the empirical formula is
c p  8.71 + 66 X 10* *  22 X 10' 1* cal mol'i deg~ l .
Find uto M75.
16. Molal heats of solids. We shall postpone the detailed theory
and discussion of specific heats till Chapter XVIII and shall give here
only two laws pertaining to solids. Though not rigorous, they are
satisfied with a fair approximation at ordinary temperatures and
are very useful because of their simplicity.
(A) The law of Dulong and Petit (1819). The molal heats of all
elementary solids have approximately the same value, close to
6 cal deg" 1 mol^ 1 . (Compare Table 5).
TABLE 5
MOLAL HEATS OF ELEMENTARY SOLID SUBSTANCES
Element
/C
C
Element
t C
C
Element
/C
c
Ag
6.00
Fe
5.85
Zn
5.99
Au
6.07
Mg
5.83
C (graph)
1.82
Bi
25
6.07
Ni
6.05
C (diam)
1.25
Cd
6.15
Pb
6.25
B
1.9
Co
5.83
Pt
6.13
Cr
5.35
Sb
6.00
Cu
5.81
Sn
6.33
(B) Neumann's law (1831). The components of solid chemical
compounds have, approximately, the same heat capacity in the
compound as in the free solid state. In other words, the molal heat,
divided by the number of atoms n in the molecule, is roughly equal to
six (Table 6).
TABLE 6
MOLAL HEATS OF SOLID COMPOUNDS
Substance
/C
c
n
c/n
Substance
/C
c
n
c/n
AgCl
28
12.53
2
6.27
PbS
12.12
2
6.06
CuO
22
10.39
2
5.20
SnS
12.63
2
6.32
CuS
25
11.89
2
5.95
ZnO
1699
10.15
2
5.08
KC1
23
12.4
2
6.2
ZnS
11.17
2
5.58
LiCl
11.96
2
5.98
CaF 2
1599
16.82
3
5.61
MnS
12.11
2
6.06
Cu 2 S
1952
19.1
3
6.37
NaBr
020
12.12
2
6.06
PbF 2
034
17.69
3
5.90
NaCl
24
12.14
2
6.07
PbCl 2
020
18.25
3
6.08
PbO
23
11.57
2
5.79
Pb0 2
24
15.50
3
5.17
46 TEXTBOOK OF THERMODYNAMICS HI 17
17. The heat function and the isobaric process. On both sides of
eq. (3.14) of JouleThomson appears the expression
X  u + pv. (3.21)
The quantity x is known as the (molal) heat function, also called heat
content and enthalpy. Like the internal energy u, the heat function is
completely determined by the variables of the system: It is another
example of a * 'characteristic function" (section 13). Differentiating
eq. (3.21), we find du = dx vdp pdv and, substituting this into
the expression (3.07) of the element of heat imparted to a simple
system,
Dq = d x  vdp. (3.22)
In a similar way, referring all quantities not to 1 mol of a substance
but to the whole simple system, we define as the total heat function
X SB JJ + pV and find DQ = dX Vdp. The heat function acquires
a particular importance in the socalled isobaric process, i.e. a process
which takes place without change of pressure (p = const, dp = 0).
For an infinitesimal isobaric change the last equation gives DQ = dX.
In the case of a finite reversible change of the system, at constant />,
from the state (1) to the state (2), we obtain by integration
Q = X 2  Xi. (3.23)
The heat imparted to the system in a reversible isobaric process is
equal to the difference of its heat functions in the final and initial states.
Under usual laboratory conditions, processes of change of the physical
state (melting, vaporization) are conducted at constant pressure, as
well as most chemical reactions. The formula (3.23) has, therefore,
a wide range of applications. For instance, the latent heat per mol is,
obviously, equal to the difference of the molal heat functions of the
material in the two states of aggregation
/ . X 2 ~ Xi. (3.24)
If we choose T and p as the variables describing simple systems,
we can write eq. (3.22) in the form
and comparing with (3.13), (3.14),
(3.26)
HI 18 CALORIC PROPERTIES OF MATTER 47
In the case of a perfect gas, the heat function can be given explicitly:
According to (1.13), x = " + RT, whence from (3.19) and (3.20)
x  / c p dT + u , (3.27)
A)
and if c p = const,
X = c p T + u . (3.28)
Exercise 10. One mol of water is vaporized at 100 C and the vapor heated to
650 C. Calculate the increase of x in this process, using the data given in exercises
15 and 17.
18. The isothermal and adiabatic processes. It will be useful to
discuss here the simplest thermodynamical processes, especially, with
respect to the amounts of heat received and work done. Since the
perfect gas is, so far, the only system whose behavior we know com
pletely, both as to its equation of state and its internal energy, we can
give explicit analytical formulas only for perfect gases. The internal
energy being a ' 'characteristic function" (compare section 13), its
change is determined by the initial (1) and final (2) state of the system
and is independent of the particular process. In the case of perfect
gases it is, according to eq. (3.18),
/ rj\ rr< \ /^ OQ\
or if the gas is only thermally perfect, and calorically imperfect,
according to (3.19),
/Tj
c,dT.
,
(3.30)
On the other hand, the work w depends on the special process and,
consequently, also the heat q which can be determined from the first
law of thermodynamics (referred to 1 mol of a chemically pure system),
q = U2 #1 + w. (3.31)
We have already discussed the isobaric process (p const) in the
preceding section. The process consisting in heating a simple system
without change of volume (V = const) is so trivial that little need be
said about it. It is called "isochoric" and, obviously, takes place
without external work: w = and q = #2 i
Of great importance is the isothermal process (T const).
We have already used the concept of isothermal curves (section 5),
regarding them as the loci geometrici of all possible states of a system
corresponding to the same temperature. However, we can interpret
48 TEXTBOOK OF THERMODYNAMICS III 18
them as the graphical representation of the following reversible process:
The system is contained in a vessel with heatconducting walls
immersed in a large heat bath of the temperature T. The geometrical
variables of the system (e.g. the volume V, if it is simple) are then
infinitely slowly changed by external forces through their whole
range of variability. If the compression or expansion takes place
with finite velocity the temperature of the system will not remain
strictly uniform, so that a process cannot be isothermal unless it
is conducted in a reversible way. In the case of a thermally perfect
gas, eqs. (3.29) and (3.30) give both u 2 u\ = 0, so that q = w.
Moreover, p = RT/v, Dw = pdv = RTdv/v, and since T is constant,
/2
= RT log   RT log  (3.32)
. V V\ p2
Equally interesting is the adiabatic process: This is the name for
anything that can happen in a system surrounded by a heatinsulating
or "adiabatic" (compare section 2) envelope which cuts off any inter
change of heat with the environment. This cover does not, in general,
prevent the system from doing work against outer forces, as is illus
trated by the example of a gas in a cylindrical vessel closed by a slid
ing piston, both of adiabatic material. The gas could do work expand
ing and lifting by its pressure a load on the piston. Unlike the isother
mal, the adiabatic process can be conducted either reversibly or
irreversibly. Since no heat is imparted in it to the system (0 = 0),
the first law of thermodynamics (3.05) takes the form
W = Ui  U*. (3.33)
The system can do work only at the expense of its internal energy.
In particular, the work of adiabatic expansion of a perfect gas from
the state (1) to the state (2) is (per 1 mol), according to (3.29),
w c v (Ti  T 2 ). (3.34)
This is the formula which Sguin was trying to verify in his attempt
to establish the first law experimentally (section 11). It implies the
same tacit assumption which was made by Mayer: he took it for
granted that the expansion of real gases proceeds without internal
work.
When the adiabatic process is conducted reversibly, to every infini
tesimal step of it applies the differential equation
Q, (3.35)
Ill 18 CALORIC PROPERTIES OF MATTER 49
and in the particular case of the compression or expansion of
a perfect gas, c v dT + Dw  0. We substitute for Dw as above, and
obtain c v dT + (RT/v)dv = 0. Dividing by c v T and using the nota
tion of section IS, 7 1 = R/c V9
y + (7~ 1)*0. (3,36)
In the derivation of this formula only the thermal equation of state
was used. It is valid also for the case where c v and 7 depend on
temperature. However, when the specific heats are constant (calori
cally perfect gas) its integral is particularlv simple. The equation is,
then, easily integrable and gives
TV 1 = const. (3.37)
We can substitute T = pv/R from the equation of state:
pv y = const. (3.38)
These two formulas are equivalent forms of the equation of the
adiabatic for perfect gases, first derived by Poisson (1823) on the
basis of the old ideas of a caloric fluid. Like the isothermals, the
adiabatics can be represented as a family of curves in the
(,p)diagram. These curves have the same asymptotes as the iso
thermals pv = const, namely the axes p = 0, v = 0; but the slope of
the adiabatic (dp/dv = yp/v) is always steeper than the slope of the
isothermal (dp/dv = p/v) passing through the same point.
One of the most accurate ways of determining the ratio 7 ** c p /c v
experimentally is the measurement of the velocity of sound in the gas
in question. The theory of elasticity gives for the velocity of sound a
the formula (already derived by Newton)
a2=s "T
M dv
where /* is the molecular weight of the gas. The rate of compression
and expansion in a sound wave is so fast that there is no time for any
appreciable heat interchange by conduction and the process must be
considered as adiabatic. 1 If the medium is a perfect gas, we find,
therefore
a 2 ypo/n. (3.40)
1 Compare section 122 with respect to the limitations of the method of sound
velocities. *4
50 TEXTBOOK OF THERMODYNAMICS HI 19
Exercise 20. Calculate numerically the work done by a perfect gas in its
isothermal expansion from i * V Q to v t  $t>o at C. Express it in ergs and in
calories.
Exercise 21. Give the expression for the work of adiabatic expansion in the
case of a Van der Waals gas. Taking the coefficients a and b from Table 2 on p. 14,
calculate it numerically for H* and N 8 and v\  v 0t v* = Jt> and obtain the difference
between the work done by these two real gases and the perfect gas.
Exercise 22. Give the general form of the integral of the eq. (3.36) of adiabatic
expansion, when c v (and c p ) is a function of T. Apply it to the empirical formula for
water vapor given in exercise 17.
Exercise 23. Calculating the work of adiabatic expansion from the formula
(3.38), check the result (3.34).
19. The cyclic process of Carnot. The French engineer S. Carnot
devised (in 1824) a process which represents a schematic approximation
of the essential features in the operation of a heat engine. Suppose
that we have two heat reservoirs of so large a capacity that their
respective temperatures Ti and T 2 (<Ti) are not appreciably changed
if we withdraw or add to them a finite amount of heat. They corre
spond in Carnot's scheme to the boiler and the cooler of a steam engine.
We shall explain Carnot's process, giving to the operating part of his
ideal engine, at first, a special construction: Let it consist of a cylin
drical vessel, with a sliding piston as its top, filled with N mols of a
perfect gas. The cylindrical side walls and the piston are made of
adiabatically insulating material while the flat bottom is heat con
ducting but can be overspread with a suitable adiabatic cover. In the
beginning the gas is in a state represented by the point A (pi, vi, T\)
of the (/>,iOdiagram (Fig. 8), and the process consists of four divisions:
(1) The bottom is heat conducting and in contact with the first heat
reservoir; by an infinitely slow reduction of the outer pressure on the
piston the gas is expanded in a reversible isothermal way at the tem
perature Ti to the point A'(p'\> v'i, Ti). (2) The bottom is made
adiabatic by applying the cover and the expansion continued, reversi
bly and adiabatically , to the point B'(p f 2, 1/2, T 2 ). (3) The bottom is
again made conducting and brought in contact with the second
reservoir; the gas is reversibly compressed along the isothermal
T T* to the state B(fa t i> 2 , T*). (4) Finally, the bottom is adia
batically covered a second time and the compression conducted,
reversibly, along the adiabatic BA. The point B is chosen in such a
way that this adiabatic brings the gas back to its initial state A, thus
completing the cycle.
The general Carnot process has in common with that just described
that it operates between the same two heat reservoirs and consists of
four analogous divisions: Two of them are an expansion and com
Ill 19 CALORIC PROPERTIES OF MATTER SI
pression of the working part of the engine, respectively, in contact
with the two heat reservoirs (boiler and cooler) ; the other two are an
adiabatic expansion and compression. However, it is less specialized
in two respects. In the first place, the working substance need not be
a perfect gas but can be any real or ideal thermodynamical system.
In the second place, the four divisions of the cycle may be conducted
not reversibly but with a finite veloc
ity (irreversibly).
Since no heat is imparted to the
working system in the adiabatic ex
pansion and compression, its total
heat intake Q during the cycle con
<P 2 'V
<P 2 ',v 2 '>
sists, in all cases, of the heat Q\
(positive) received from the reservoir
of higher temperature (T\) and the v
heat Q 2 (negative) received from that FIG. 8. Carnot's cyclic process.
of lower (r 2 ), so that Q = Qi + Q 2 .
As the final state of the working system is identical with the initial,
we can apply the equation of the first law in the form (3.06) :
W = Q = Qi + Q 2 .
From the standpoint of the engineer, Qi is the heat conveyed to
the engine to make it operate and supplied by the consumption of fuel.
Not all of it is transformed into work, because Q 2 is negative. The
heat given to the cooler ( Qz) is, to his mind, a wasteful loss of energy
which is, however, necessary in order to bring the working part of the
engine back to its initial state and to ensure a continuous operation by,
periodically, repeating the cycle. The ratio rj of the useful work W to
the consumed heat Qi is called the efficiency or the conversion factor
of the engine
W Qi +
We shall compute the efficiency for the special case when the cyclic
process is conducted reversibly with N mols of a perfect gas as operat
ing substance (Fig. 8). From the expression (3.32) of the heat imparted
to the gas in an isothermal process
Q l  NRTi log (v'M, Q 2   NRT 2 log (v' 2 /v 2 ).
On the other hand eq. (3.37) applied to one of the adiabatic divi
sions gives 7W" 1 * T 2 v 2 y " 1 or (vi/vtf* 1  T 2 /Ti, and to the
52 TEXTBOOK OF THERMODYNAMICS III 19
other: (i/i/fl^) 7 " 1 = T*/T\. There follows vi/v* = v'i/v' 2 or
v'i/vi = v'z/V2. Consequently
Qi + Q2 = NR(Ti  T 2
and
QI + & Tt T 2
V = ^  * ^ , (3.42)
Qi Ti
a relation which can, also, be written in the form
+ 0 (343)
1 1 12
and will prove very important in the next chapter.
Exercise 24. Use the formula of adiabatic expansion derived in exercise 22 to
show that the relation v'\/v\ v'z/v^ holds also in that case. The efficiency of
Carnot's process is the same when c v is a function of T.
Exercise 25. What is the thermodynamical efficiency of engines having coolers
at C and boilers at 100 C, 150 C, 400 C?
CHAPTER IV
THE SECOND LAW OF THERMODYNAMICS
20. Several formulations of the second law. The second law of
thermodynamics is older than the first. It was discovered by S. Car not,
who enunciated it in the same pamphlet (1824) in which he gave the
theory of his cyclic process. The reinterpretation and extension of
the second law on the basis of the energy concept is due to Lord Kelvin
(1851) and Clausius (18501863). There exist several equivalent for
mulations of this principle of which we give here the following:
1 st enunciation (Kelvin) : It is impossible to transfer heat from a
colder system to a warmer without other simultaneous changes occurring
in the two systems or in their environment.
The reader will recognize that this is a more precise and amplified
statement of one of the facts which we mentioned in the very beginning
of this course (section 1), where we introduced the notions of equi
librium and temperature, and again in section 10, where we defined
the concept of heat. We said there that heat flows spontaneously
always from the higher temperature to the lower and never in the
opposite direction. The second law goes beyond this and asserts that
one cannot completely undo the results of the process of heat conduc
tion even in any indirect way. We shall show in the following sections
that the above formulation is completely equivalent to the
2 nd enunciation : // is impossible to take heat from a system and
to convert it into work without other simultaneous changes occurring in
the system or in its environment.
This law does not prohibit a process like the adiabatic expansion
of a system (in which work is done at the expense of heat in the form of
its internal energy) because this process is attended by " another simul
taneous change", namely, the change of volume of the system. Neither
does it forbid the operation of Carnot's engine (preceding section):
In his cycle heat is taken from one reservoir, but its conversion into
work is not complete. The simultaneous change, which makes this
process permissible, is the transfer of part of the heat to another
reservoir of lower temperature. The transferred heat is less available
for conversion into work in a similar way because this would require a
63
54 TEXTBOOK OF THERMODYNAMICS IV 21
third heat reservoir of still lower temperature. It is, therefore, often
said that this heat is "dissipated" or "degraded". In general, a
difference of pressures or of temperatures is necessary to obtain work;
therefore, the phenomena of internal friction and of heat conduction
which tend to reduce or destroy these differences must be regarded as
attended by dissipation of energy. The establishment of the first law
of thermodynamics made an end to all hopes of building a perpetual
motion machine (of the first kind) which could continually do work
without some other form of energy being supplied to it. However, for
all practical purposes a machine which could convert into worjc the
enormous stores of internal energy contained in the oceans, the atmos
phere, and the body of our earth would serve just as well. Ostwald
called such a machine, which could work without temperature dif
ferences, a perpetualmotion machine of the second kind. It is obvious
that its operation would be in violation of the second law, so that we
arrive at the
3 d enunciation (Ostwald) : // is impossible to construct a perpetual
motion machine of the second kind.
This formulation makes the second law analogous to the first, which
asserts the impossibility of a perpetualmotion machine of the first kind.
41. Efficiency of heat engines. Equivalence of the formulations.
We have calculated, in section 19, the efficiency y = (Q\ + (?2)/(?i of
the Carnot engine when operated reversibly with a perfect gas as the
working substance. We shall call this particular type of construction
and operation the perfect Carnot engine. The second law of thermo
dynamics permits us to say something about the conversion factor of
the general Carnot cycle, which is the schematic prototype of any heat
engine. In the real case, when the working part is any thermodynami
cal system and the cycle conducted either reversibly or irreversibly, we
shall denote the efficiency by vj' = (Q'\ + Q'd/Q'i Let us suppose,
for a moment, that 17' > t?. We shall prove that this supposition is
impossible, by carrying through the cyclic process alternately with a
real and a perfect Carnot engine which share the same two heat reser
voirs. The length of the isothermal divisions A A' and B'B (Fig. 8)
of the perfect engine is at our disposal and can be adjusted in such a
way as to make the work done by the two cycles equal :
W W or Q'i + Q't  Qi + Q 2 . (4.01)
Moreover, the perfect engine being reversible, it can be run back
ward, in which case the quantities Q\, (>2 W simply reverse their signs,
so that work is consumed and heat produced in the process. What
IV 21 THE SECOND LAW OF THERMODYNAMICS 55
will now be the result of a double cyclic process in which the real
engine goes through one cycle forward and the perfect backward? No
work will appear in the net balance, because the first engine does as
much work as the second consumes. The only effect will be that the
heat Q'\ Q\ will be taken from the first reservoir and Q f 2 Qz
from the second. These quantities are oppositely equal, according to
eq. (4.01): Q f \ Qi = (Q f 2 (?2). Let us now recall our supposi
tion if > TJ or (Q'i + Q f 2)/Q'i > (Qi + (?L>)/(?I. Since the numerators
are equal, this leads to 1/Q'i > l/(?i or Q'\ < Q\ and, consequently,
Q f 2 > (?2 A positive amount of heat Q'% Q% is taken from the
reservoir of lower temperature T% and transferred to that of higher Ti
without any other changes occurring, either in the reservoirs or in
the working parts of the engines. The supposition TJ' > 17 leads,
therefore, to a contradiction with the second law of thermodynamics,
and we must conclude that it is impossible and that only the possibilities
*' ^ * (4.02)
are permissible.
In the particular case, when the process in the real engine is also
conducted in a reversible way, we can run the perfect engine forward
and the real backward. Repeating the same argument with transposed
(primed and unprimed) symbols, we can, then, prove that 17' < 17
is impossible and that 77' S must hold. This result is compatible
with (4.02) only when we have the sign of equality: rf = 77. All
reversible Carnot engines have the same efficiency, no matter what the
construction of their working part is. The sign < in eq. (4.02) refers,
on the contrary, to irreversible cycles. Summarizing these conclusions
and making use of the expression (3.42), we can give the general law
for the efficiency 17 of any engine carrying out the cyclic process of
Carnot:
(4.03)
where the sign of inequality (<) refers to irreversible, that of equality
{ = ) to reversible, operation. In the reversible case, there holds also
the relation
+ % = 0, (4.04)
ll 12
whicb it is only another way of writing eq. (3.42).
An irreversible Carnot process is always less efficient^rfhQMaocr'
wasteful than a reversible inasmuch as a smaller fraction/^ tHB Be*f (?
56 TEXTBOOK OF THERMODYNAMICS IV 22
taken from the first reservoir (boiler) is converted into useful work.
However, eq. (4.03) with the sign < will also apply when only one
of the four divisions of the cycle is conducted irreversibly and the
three others reversibly. We can conclude from this that in the expan
sion of a system in contact with a heat reservoir less heat is received
by the system and less work is done when the expansion is conducted
in an irreversible than in a reversible way between the same initial
and final states. In fact, let these two kinds of expansion be the first
divisions of two Carnot cycles, while the remaining three divisions are
identical and reversible in both. We denote the heat items in the
first (partially irreversible) cycle Q'\, Q% and in the second Q\, Q%.
The condition 17' < 77 or (Q'i + Qd/Q'i < (Qi + Qz)/Qi leads to
Q'i < Qi, since 2 is negative. The change of internal energy is in
both cases the same, because it depends only on the initial and final
state, so that the smaller heat intake is attended also by the smaller
production of work.
We have used above the first enunciation of the second law given
in section 20. However, the second leads to, precisely, the same
results. Their equivalence is shown, in a general way; by the following
simple reasoning. Consider a process which contradicts the second
enunciation, that is, one which produces the work W taking heat from
a reservoir at the temperature T^ without any other simultaneous
changes. We can feed this work into a reversible Carnot engine
which runs in the backward direction, adjusting it so that in a cycle
just the work W is consumed, and using the reservoir T<z as the cooler.
The net result will be that a certain amount of heat will be transferred
from this reservoir to one of higher temperature TI (the heater of the
Carnot engine), ^'without any other simultaneous changes". In
other words, a process which violates the second law in the sense of
enunciation 2, violates it also in the sense of enunciation 1. As to the
third enunciation, it is in substance identical with the second being,
merely, a restatement of it in different words.
Exercise 26. Give an alternative proof of the relation (4.02), for the second
enunciation, analogous to that given in the text, for the first. Instead of adjusting
the two engines so as to satisfy (4.01), adjust them so that the total heat received
by the cooler in the double cycle vanish.
22. The concept of entropy. All that follows from the second law
of thermodynamics for a reversible Carnot cycle is embodied in the
formula (4.04). We now extend our considerations by asking: What
follows from it in the case of any other reversible cyclic process (different
from that of Carnot's) carried out with a thermodynamical system.
Our arguments and formulas will refer to systems of the most general
IV 22
THE SECOND LAW OF THERMODYNAMICS
57
kind, but we shall illustrate them by a drawing (Fig. 9) relating to a
simple system, the only one which can be represented in a two
dimensional diagram. We repeat what was said in section 13 about
carrying out such a process. We must have an infinity of heat reser
voirs differing in temperature by infinitesimal amounts. The system
is brought in contact with them, successively, and at the same time
subjected to infinitely slow compressions or expansions. The cycle is,
then, visualized by the solid closed curve of the figure. In its infinites
imal part corresponding to the segment A A' the system receives an
element of heat DQ. Let us now draw through the point A the isother
mal AC and through the point A' the adiabatic A'C, and denote by
DQi the heat which the system would receive if it were subjected to
the infinitesimal isothermal process AC.
The relation between DQ and DQ\
is obtained by considering the little
cyclic process AA'CA and setting up
the equation (3.06) of the first law
for it: the isothermal part AC is
carried out in the reverse direction and
we have DQ  DQi = DW. The ex
pressions DQ and DQ\ are infinitesimal
of the first order, but the work DW
done in the little cycle is represented
by the area of the triangle AA'C and is,
therefore, infinitesimal of the second
order, infinitely small compared with
the other two terms. It must be omitted from the equation, leaving
DQ = DQ\: we can, therefore, change our notations, denoting the
work done in the segment A A' also by DQ\. This conclusion holds
for any system and not only for a simple one: Although the work
done in an infinitesimal cyclic process cannot, in general, be re
presented in a twodimensional diagram, it is, nevertheless, always
infinitely small of the second order.
Let us now continue the adiabatic through the point A 1 and draw
one through A until they intersect the closed curve a second time in
B', B. According to the result just obtained, the heat DQ^ received
by the system in the part of the process B'B is the same which corre
sponds to the isothermal segment B'D passing through B f . We see,
therefore, that the two heat items DQi and DQ* are precisely the same
which the system would receive from the boiler and cooler if it formed
the operating part of a Carnot engine carrying out the reversible Carnot
cycle ACB'DA. Denoting by T\ and T* the temperatures which the
\ B
FIG. 9. General cyclic process.
58 TEXTBOOK OF THERMODYNAMICS IV 22
system has in the points A and B', we can apply to DQ\ and DQ2
the relation (4.04) for a reversible cycle of Carnot
(4.05)
JL I J. 2
If we cut up into strips the whole area of the closed curve repre
senting the cycle, by a system of adiabatics drawn at infinitesimal
distances from one another (the dotted lines of Fig. 9), eq. (4.05) will
hold for the pair of segments of the cycle in every strip. Integrating
over all strips
C no* C nn*
= o.
The first integral represents the summation over all elements of
the cycle between the points M and N, the second term the summa
tion over the other half of the curve from N to M. We can, therefore,
replace the two terms by the single integral
0, (4.06)
extended over the whole cycle. This result applies to any cyclic
process, so that the integral of the linear differential expression DQ/T
taken over any closed path vanishes. We know from section 8 that,
in this case, the expression DQ/T is the exact differential of some
function 5 of the variables of the system :
DQ dU + DW J0
j   J>   dS. (4.07)
In the language of that section we say that DQ, while not an exact
differential, is integrable. The integrating multiplier is 1/T, the
reciprocal of the absolute temperature as measured by a perfect gas
thermometer.
Clausius who first introduced the function 5 called it the entropy
of the system (derived from the Greek word Ivrpcnofuu = turn inside,
an allusion to its onesided character). If we integrate eq. (4.07) over
a reversible process leading from state (1) to state (2), we find
 Si  . (4.08)
J\ T
The integral depends only on the initial and final value and is
independent of the path, provided that this path is reversible (compare
IV 22 THE SECOND LAW OF THERMODYNAMICS 59
section 8), and the entropy is a "characteristic function", i.e. com
pletely defined when the state of the system is defined. We can take
as state (2) in formula (4.08) any state of the system and as (1) some
state (n) which we arbitrarily choose as normal
(4.09)
(n)
Owing to the fact that the entropy is defined by the differential
equation (4.07), it contains a constant of integration. As such ap
pears in the expression (4.09) the entropy S n of the normal state. As
.far as the first and second laws of thermodynamics are concerned, it
cannot be determined by any theoretical reasoning.
In the case of a homogeneous system, the heat DQ received in
every element of a reversible process is distributed uniformly, so that
the entropy S also acquires a uniform distribution over the system.
We can speak, therefore, of the specific entropy and, when the system
is chemically pure, of the entropy per mol or molal entropy which we
shall denote by 5 (small). We may further ask: What is the joint
entropy S of two systems which, individually, possess the entropies
Si and 52? According to the defining equation (4.09), S is obtained
by carrying out a reversible process with the combined system. The
reversibility requires, however, the equilibrium of all parts at every
stage of the process, so that our two constituent systems must have
at every moment the same temperature T. If we denote the elements
of heat imparted in an infinitesimal step of the process, respectively,
by DQ, DQi, DQ 2 , there follows DQ  DQi + DQ 2 , DQ/T  DQi/T
+DQz/T and, from eq. (4.07),
dS = dSi + dS 2 , or S = Si + S 2 . (4.10)
The entropy is additive. We repeat, however, that the additive
property rests on the fact that two systems in thermodynamical
equilibrium have the same temperature. If two systems are, initially,
not at the same temperature but can be brought into thermal contact,
it has a good sense to ask about their joint entropy, and we shall define
it as additive. In other cases this question is without meaning. It is,
of course, possible to imagine a case in which the two parts are adia
batically insulated from one another but can interact by means of
movable pistons and similar devices. They could be in partial equi
librium without equality of temperature, but this is a dynamical
problem rather than a thermodynamical. The notion of an entropy
TEXTBOOK OF THERMODYNAMICS
IV 23
of such a combination is artificial and without interest in thermo
dynamics. 1
The entropy difference (4.08) takes a very simple form when the
initial and the final states are of the same temperature. We can, then,
take an isothermal process as the path of integration (T = const).
S2 Si = 7T
(4.11)
(S r T 3 )
23. The entropy principle. Let us discuss how the entropy of a
system changes in an adiabatic process. If the process is reversible,
eq. (3.35) or DQ = applies to every infinitesimal step of it. It
follows, then, from (4.07)
dS = 0, and S = const. (4.12)
The reversible adiabatic process leaves the entropy unchanged: it is,
therefore, also called " isentropic".
When we turn to the irreversible adiabatic process, we are on new
ground. We cannot expect to get an answer to our question from the
considerations of the preceding sec
tion because they were restricted to
reversible processes. Therefore, we
have to fall back, once more, upon
the second law as enunciated in
section 20. In the same way as in
introducing the entropy concept we
shall use a (p, V) diagram (Fig. 10)
as an illustration, although our ar
gument will be valid for the most
general thermodynamical systems.
Suppose that the system is sub
jected to an< irreversible adiabatic
process which brings it from the
state A (with the temperature T\ and entropy Si) to the state
B(T%, ^2). The transition itself, being irreversible, is not capable of a
graphical representation (compare section 12). What can be said
about the entropy change AS = 2 ~ Si? We can bring the system
1 The only known significant case in which the entropy is not additive occurs in
the theory of radiation. When a pencil of rays is split by partial reflection into two
(the reflected and retracted), they have, in general, temperatures different from each
other's and from that of the parent pencil. At the same time the sum of their
entropies is not equal to the entropy of the parent pencil. Compare: M. Von Laue,
Ann. der Phys. 20, p. 365 f 1906; P. S. Epstein, Phys. Zs. 15, p. 673, 1914.
FIG. 10. Entropy change in irre
versible adiabatic process.
IV 23 THE SECOND LAW OF THERMODYNAMICS 61
back into its initial state A in the following reversible way: First,
we let it undergo a reversible adiabatic process EC roughly opposite
to the first, until it comes to the temperature Ta, while its entropy
remains constant and equal to 52. From the state C on, we conduct
it in a reversible isothermal (CD) until the initial entropy Si is again
reached; then we complete the cycle, bringing the system adiabatically
back to the state A. The temperature T% can be chosen arbitrarily
higher or lower than either T\ or 7^2, or between these values. It is
important to have this freedom of choice because certain processes
(as transformations) can be conducted only at special temperatures.
The system receives no heat in the three adiabatic branches of the
cycle, and the heat Q in the isothermal part. Applying eq. (3.06) of
the first law to the cyclic process (from A to A) we write
where W is the total work done by the system. If Q and W were
positive, the result of the process would be that the heat Q is taken
from the heat bath, maintaining the constant temperature T% in the
isothermal branch CD, and converted into work " without any other
simultaneous changes." This is a violation of the second law (enunci
ation 2), and we conclude Q > or Q ^ 0. The application of
eq. (4.11) to the isothermal CD gives Q = Tz(Si 2), since Si is
here the final and 2 the initial state. The inequality implies, therefore,
Si  S 2 ^ or
AS = S 2  Si ^ 0.
The same argument can be easily adapted to the more general case
where the system consists, initially, of many parts which are not in
equilibrium. At the end of the irreversible adiabatic process it may
have the same parts in new states or, even, other parts since some of
the parts may have become merged, others separated into several units.
At any event, the system can be brought back to the initial condition
by the following reversible procedure. The final parts are brought,
separately, in a reversible adiabatic way to the same suitably chosen
temperature TV At this temperature they are iso thermally rearranged
into the original parts, unmixing some of them if necessary (compare
section SO). Still at the same temperature T$ each of the part systems
is separately brought to its initial entropy. The cycle is completed by
bringing the parts back to their initial states in independent, reversible,
adiabatic processes. The conclusions drawn above for the simpler
case obviously apply without change in the more complicated one.
They apply, of course, also to every part of the irreversible process for
which a definite entropy change can be defined. Under no circum
62 TEXTBOOK OF THERMODYNAMICS IV 23
stances can this change be negative. It follows from this that, in the
case 52 Si = 0, the change must be zero for every smallest part:
dS = 0. That is to say, A5 = refers to the reversible process and
A5 > to the irreversible : The irreversible adiabatic process is always
attended by an increase of entropy.
The summary of the results of this and the preceding section is the
"entropy principle", due to Clausius, which represents an analytical
formulation of the second law of thermodynamics. It consists of two
parts:
4 th enunciation: (a) There exists a characteristic thermodynamical
function called the entropy. The difference of entropies of a system in
the states (1) and (2) is given by the expression
(4.13)
where the integral is to be taken over any reversible path connecting the
two states.
(b) In an adiabatic process the entropy either increases or remains
unchanged:
AS ^ (4.14)
where the upper sign (>) refers to the irreversible, the lower ( = ) to
the reversible, case. The adiabatic processes include also the case when
the system is "left to itself", meaning by this that it is completely
cut off from any interaction with its surroundings and does not receive
any energy from outside, either in the form of heat or of work.
The parts (a) and (b) of the entropy principle contain all that can
be deduced from the enunciations of section 20. Sometimes, a third
general law is included in the entropy principle: the fact that all
systems tend towards equilibrium so that the entropy always does
increase when the conditions of the system permit it. We prefer, how
ever, to discuss this law separately in section 30. As enunciated here,
the entropy principle is completely equivalent to the formulations
given in section 20. In fact, their validity is a sufficient condition
because we derived the principle from it. But it is also a necessary
condition because the adiabatic processes prohibited in those formu
lations are attended by a decrease of entropy. (Compare exercise 27).
Since a spontaneous irreversible process is one of increasing entropy,
we can use the condition AS > as the criterion of irreversibility.
Exercise 27. Show that, in the fictitious case when the efficiency of the Carnot
engine is y f > ij, the expression on the left side of (3.43) is negative. This means
that the sum of the entropies of the heater and cooler decreases.
IV 24 THE SECOND LAW OF THERMODYNAMICS 63
24. Entropy of the perfect gas. In the case of a simple homo
geneous and chemically pure system, the entropy can be referred to 1
mol of its material (section 23), and eq. (4.07) takes the form
(4 , s)
The only system for which the expressions of u and p were given
in the preceding chapters is the perfect gas: According to (1.13) and
(3.17) we have du = c v dT and p = RT/v t
<fe*y + ** (4.16)
This, really, is an exact differential of the function
s = c v log T + R log v + s'o, (4.17)
where s'o is an arbitrary constant, the socalled entropy constant of the
perfect gas. Using the equation of state, we can express v by p and T,
or T by p and v, and so obtain two other forms of the molai entropy
s = c p log r  R log p + so, (4.18)
s = c v log p + c p log v + s"o, (4.19)
since c p = c v + R, according to (3.20). The constants are related
in the following way: s'o = so R log R, s"o = so c p log R. It is
easy to see that eqs. (3.37) and (3.38) of the adiabatic are equivalent
to the condition 5 = constant.
It is well to mention also the case when the gas is perfect, in thermal
respects, but the molal heat c v is a function of temperature (end of
section 15). Equation (4.16) leads then to the expressions
(4.20)
So.
fcJT
J ~T~~
Exercise 28. Check the statement that the process of expansion into the void
(section 15) from the molal volume Vi to v 2 is irreversible (i.e. As > 0) for the perfect
gas, by computing the entropy difference $ 2 SL between the final and the initial
state.
Exercise 29. Two vessels contain each one mol of the same perfect gas at the
respective temperatures Ti and TV They are brought in contact and the two
specimens allowed to come into equilibrium through the heatconducting walls
without change of volume. Compute total entropy in initial and final states and
64 TEXTBOOK OF THERMODYNAMICS IV 25
show that it increases. (Remark: The arithmetical mean of two numbers J(a + b)
is always larger than the geometrical *
25. The reciprocity relation of thermodynamics. The require
ment that the expression (4.15) must be an exact differential imposes a
restriction on the analytical form of the internal energy u: the caloric
equation of state (3.09) is not quite independent of the thermal. In
fact, the exact differential must satisfy the reciprocity relations (2.10)
of section 8. In the case of a simple system, there are only two inde
pendent variables and only one reciprocity relation. Let us substitute
for Dq the form (3.11)
*?()
3 T. dT +  n ^
Then the reciprocity relation has the form
~ir\() + T\' (4 ' 22)
: oi LJ. \OV'T 1 Jt>
In carrying out the differentiation, we shall remember, that u is
here regarded as a function of v and T only, and drop the subscripts
in the second derivative
1^2 1
O ** 1 ^, ^.
+
l(*P\\(*\ + J
T\dTj v T 2 l\dv/T *T
The order of differentiation in the second partial can be transposed
and the two terms depending on it cancel out. There remains
.*
which gives, substituted into (4.21),
(IX ();
In the case of perfect gases, (dP/dT) v = p/T and (du/3v) T = 0.
We justified the assumption c v = f(T), for them, by the measurements
of Joule and Thomson on real gases (section IS). We see that this
assumption is consistent with the second law, as we have already
checked in the preceding section. Vice versa, accepting the second law,
we can conclude theoretically that u is independent of v and a function
of T only, for substances satisfying the equation pv = RT. This fact
is well supported by observations on real permanent gases. When
IV 25
THE SECOND LAW OF THERMODYNAMICS
65
the measurements are very accurate, the deviations from the law of
perfect gases begin to tell. The incentive which caused Thomson and
Joule to undertake their investigation was the fact that Regnault's
data for (dp/dT) v , when substituted into (4.23), gave a finite value of
(du/dv)r for real gases.
In the case of gases strictly obeying the eq. (1.21) of Van der
Waals, we find from (4.23)
= J' < 4 ' 25 )
This partial is independent of T. Since du is also an exact differen
tial, we can apply the reciprocity relation to it and conclude that the
other partial c v = (du/QT) v is independent of v. The internal energy
has then the expression
/
c v dT  
(4.26)
and the term a/v represents the inner potential energy of the forces
of interaction between the molecules.
Equation (4.23) can be used to simplify the expression (3.15) of
the difference of molal heats c p c v , making it dependent upon easily
measurable quantities
This formula was extensively tested by measurements o/ the
velocity of sound in gases (section 18) up to pressures of 200 atm and
found to be in good agreement with observations. 1
Exercise 30. Show that the entropy of the Van der Waals gas is
(4.28)
Exercise 31. Derive by integrating (4.23)
Exercise 32. Starting from (4.23) prove the relation
P. P. Koch, Ann. Phys. 26, p. 551; 27, p. 311, 1908.
66 TEXTBOOK OF THERMODYNAMICS IV 26
Exercise 3d. A further relation can be derived by partial differentiation. Sub
stitute in u( T, v) the expression of T in terms of p and v so that u = u[T(p t v) t v].
By differentiating with respect to v derive
(Q\ _ (^u\ ( ^ u \ (^T\
9v/ p \ Qv/7 1 \97/ v \9i/ /p
Transform this with the help of (3.12), (4.23), and (4.27) into
 P (4.31)
26. Measurement of entropy of differences. From the definitions
of molal heat (3.08) and of entropy (4.15) there follows the relation
c = Tds/dT, which is valid for all ways of imparting heat to the
system. In the two simplest cases of constant volume and of constant
pressure this leads to
<4  32 >
Dividing by T and integrating, we obtain the following two expres
sions for the entropy
(4.33)
P), (4.34)
whose equivalence with (4.32) can be checked by differentiating them
partially with respect to T, at constant v and p, respectively.
If we wish to find the entropy increase (per mol) experienced by a
substance when its temperature is raised, at constant volume, from T\
to Tg, the first equation gives
/TZ
T
I
s(r a , r)  s(Ti, v) = / ^dT. (4.35)
*/ Tl 1
On the other hand, the increase at constant pressure is from the
second equation
/T2
%

, p)  s(ri, p) = / ~ dT. (4.36)
*/ TI 1
The above expressions give the entropy changes, as long as the
system does not undergo any transformation of state. When a trans
formation occurs, it involves a latent heat / (per mol) and an increase
of entropy which is, according to (4.11),
**i, (4.35')
IV 26 THE SECOND LAW OF THERMODYNAMICS 67
since the transformation takes always place at constant temperature.
Here, too, the cases v = const and p = const must be distinguished.
At constant volume no work is done and the latent heat is, simply,
equal to the increase of the internal energy. At constant pressure it
is equal to the increase of the heat function (compare section 17):
/ = Au, l p = A X . (4.360
We see from this discussion that the determination of entropy dif
ferences involves only measurements of molal and of latent heats.
The same data which are necessary to find internal energies and heat
contents give also the knowledge of entropies.
We have tacitly assumed that the system we are dealing with is a
chemically pure substance. If this is not the case, we can readily
adapt eqs. (4.35) and (4.36) to any system in which no chemical
reaction takes place, simply writing the total entropies and heat
capacities (capital letters) instead of the molal. With respect to
chemical reactions, two cases may occur: (1) The reaction is conducted
at constant temperature. This case is completely analogous to that
of a transformation, the entropy increase being
S 2 Si = , (4.35")
with
Q v = A 7, Q p = AX. (4.36")
(2) There is a continuous displacement of chemical equilibrium as
the temperature is raised. Then the heat of reaction must be included
in the heat capacity; we shall consider a case of this sort in section 121.
Exercise 34. Calculate the entropy increase in the following process: 1 mol
of liquid water (18 g) at 100 C is vaporized at the pressure of 1 atm. and the vapor
is heated (at constant pressure) to 650 C. Take data from exercises 8 and 19.
Exercise 35. (a) The molal heats c p for Ag, Al, and graphite are:
rK
50
100
150
200
250
273.1
298.1
Ag
2.69
4.82
5.54
5.84
5.97
6.02
6.04
Al . .
90
3.13
4.44
5.13
5.54
5.68
5.831
C (graphite) . .
. 0.13
0.41
0.79
1.22
1.65
1.86
2. fed
Calculate in a rough graphical way the increase of the heat function as
temperature is raised from 50 to 298.l K (p = const). Ot the
(b) Calculate from the same data (in a rough graphical way) the in'
entropy between 50 and 298 .l K.
Exercise 36. Calculate the heat and the entropy imparted to a v~** (4.43)
the following two reversible processes: (a) starting from the state pj
is conducted first at p PI to pi, , then at v vt to pi, PI. (V 'The introduc
68 TEXTBOOK OF THERMODYNAMICS IV 27
at v v\ to pz, Vi, then at p pi to pz t vz. Suppose c v and c p to be constant.
Show that the heat imparted to the system is different in the two processes but the
entropy change is the same.
27. The extended law of corresponding states. The relation
(4.23) between the caloric and the thermal equations of state can be
written in the integrated form
= f [ T (^  p]dv + f(T).
(4.37)
We have seen in section 6 that, within certain groups of substances,
the thermal equation satisfies the law of corresponding states. It may
be asked, therefore, if a similar law does not apply to the caloric
equation of state (4.37). As the most precise expression of the law of
correspondence we may take that following from the mean reduced
equation (1.29) of Kamerlingh Onnes. According to that equation, the
quantity K RT c /p c v c is a constant for a given substance and, at the
same time, a universal function of the reduced variables ?r, <p, r. Conse
quently, it must be a universal constant, as also
^ = ., (4.38)
L c
since K = R/K is another universal constant. If we substitute into
the combination pv/T the expressions of the parameters in terms of
the reduced variables (p = p c TT, v = v c <f>, T = T c r), we find im
mediately
^ = K^ = /i(ir,*>, r). (4.39)
In words, the quantity pv/T has the same value for different substances
in corresponding states. In a similar way, substituting TT, #>, r into
the integral of eq. (4.37), we find that this integral, divided by T,
is a universal function of the reduced variables. This fact suggests
he hypothesis that the same applies to the second term and to the
>ole expression (apart from the zero energy)
jT =/2(7r,^r). (4.40)
iantity might be also the same for different substances in
states. This hypothesis was, in fact, made by
nnes and others. We shall call the simultaneous validity
IV 28 THE SECOND LAW OF THERMODYNAMICS 69
of eqs. (4.39) and (4.40) the extended law of corresponding states. Though
far from rigorous, this law is satisfied with a fair approximation within
certain ranges of temperatures by large groups of substances (compare
section 47). It may be pointed out that, in the case of perfect gases,
the two functions /i and /a have the simplest possible form : they
are reduced to the constants R and c v .
An immediate consequence of these equations is that the expres
sion (4.15) of the entropy differential is also a universal function Fof
TT, <p, T, whence
As = Ffa, <p2j r 2 )  F(TTI, ?i, n). (4.41)
The molal entropy difference between corresponding states is the same
for all substances (obeying the extended law).
There were physicists who went even beyond this in their extension
of the law and applied it to more general systems, having other
variables than />, V, T. Suppose, for instance, that the element of
work is, in the sense of section 7, dW = pdV + ydX, whence the
entropy differential (4.07)
dU + pdV + ydX
dS =    (4.42)
The generalized force y can be treated exactly on the same footing
as p (exercise 38), and it is possible to argue that the expression yX/T
should have similar properties (with respect to corresponding states)
as pV IT and U/T. We shall see in sectioit 86 that this idea was
fruitfully applied to the theory of surface tensions.
Exercise 37. Substitute for p, v, T the reduced variables into the integral of
eq. (4.37), divided by T, and show that the result is a universal function of IT, <p, r.
Exercise 38. Substitute in (4.42)
v,x /x,T \dX/ T .v
and set up the reciprocity relations in analogy with (4.22), (4.23).
28. Theory of the JouleThomson process. We are now in a
position to subject the experiments of Joule and Thomson, described
in section IS, to a closer analysis. We found there that the relation
between the states of the streaming gas (Fig. 7) on both sides of the
plug is given by eq. (3.16), which we can write in the form
AX = X2 xi = 0, (4.43)
making use of the definition (3.21) of the heat function. The introduc
70 TEXTBOOK OF THERMODYNAMICS IV 28
tion of this function permits us to express the entropy differential in
a new way: From eqs. (3.25) and (4.15)
We apply the reciprocity condition (2.10) to this expression and
obtain, in complete analogy with the way in which we found the for
mulas (4.23) and (4.15),
, .
and
Of course, this is not a new relation but only another form of the
reciprocity condition (4.23) into which it can be transformed by
applying the formulas (1.05) and (3.21). Owing to this equation and
to (3.26), the differential of the heat function becomes
[v  T (j) ]dp.
d x = cJT + v  T dp. (4.47)
Suppose that in the JouleThomson experiment the difference of
pressures A/> = p2 pi and of temperatures AT 1 = T% T\ is so
small that it is permissible to neglect squares and higher powers and
use eq. (4.47) for expressing AX in terms of &p and A7\ The condition
(4.43) gives then
Ar v  T(Qv/dT) p
" = ^ =  7 P  ' (4 ' 48)
This change of temperature per unit pressure drop, n = AjT/A/>,
is, usually, called the differential JouleThomson effect. With the help
of this formula its sign and magnitude can be calculated theoretically
if the equation of state and c p are known. We see that the sign of the
effect depends on that of T(dv/dT) p v, or, transforming this by
means of (1.05), upon the sign of the inequality
The differential JouleThomson effect has the sign opposite to that
IV 28 THE SECOND LAW OF THERMODYNAMICS 71
of the bracket expression (because dp/dv is negative). In the case of
Van der Waals gases, this condition takes the form
2a RTb > ^
2 2 < 
v 2 (v b) 2 <
When the density is low, the constant b in the denominator of the
second term gives rise to corrections of the second order only, and can
be neglected in the first approximation, leaving the condition
T T^ Ti = 2a/Rb = 6.757;, (4.51)
taking into account (1.24). Ti is called the inversion temperature:
below it, the effect is positive and consists in a cooling of the gas;
above, it is negative and the gas gets warmer in passing through the
cotton plug. We see from (4.50) that the two Van der Waals constants
have an opposite influence, so that the sign depends on which of them
dominates. From the physical point of view, we have here an interplay
of two phenomena: On the one hand, when a is different from zero,
the expanding gas does work against inner forces (compare eq. (4.26)
of preceding section) and this decreases its temperature accounting for
a part (a/v 2 ) of the expression (4.50). On the other hand, the outer
work p22 piv\ of the process does not vanish for Van der Waals
gases, even when the temperatures remain unchanged. It can be
positive or negative and contributes the remainder of (4.50). On the
whole; the larger part of a positive effect is due to inner work.
The interplay of the two effects is more complicated when the
density is not low. The expression (4.50) with the sign of equality,
determining the inversion point, is best written in terms of the reduced
variables (1.25)
 < 4  52 '
Eliminating <f> from this expression and the reduced Van der
Waals equation (1.26), we find
(4.53)
This formula shows us that there are two inversion points. The
coefficient /* is negative above the upper inversion point and below the
lower, being positive in between. The position of these points depends
on the pressure : the upper of them sinks as the pressure is increased,
the lower rises.
72
TEXTBOOK OF THERMODYNAMICS
IV 28
In a qualitative way these predictions of the Van der Waals theory
are fulfilled by real gases. The lower inversion lies in the liquid state
and has been observed in several substances. It is, further, borne out
by the experimental data that, the more perfect a gas is, the lower its
upper inversion point lies and that it is depressed by pressure, while
the lower point is raised by it. However, the Van der Waals equation
cannot be used for a quantitative calculation of the JouleThomson
effect or of the inversion point because its approximation is insufficient
in the range of pressures and temperatures here involved. This is
apparent from the figures given in Table 7.
TABLE 7
DIFFERENTIAL INVERSION POINTS
Substance
p aim
TiK
Authority
Upper
He
1
23.6
Roebuck and Osterberg 1934
H 2
113
(192.7)
Olszewski 1901
Air
150
553
Hoxton 1919
< t
203
523
Lower
A
250
164
Roebuck and Osterberg 1933
4 4
50
125
Air
150
140
Hausen 1926
4 I
125
133
CO,
18 to 100
249
Burnett 1923
Hoxton 1 gives the following empirical formula for air, in the range
from to 280 C and from 1 to 220 atm,
(4.54)
1 J. ~
where p is measured in atmospheres.
The cooling by means of the JouleThomson process has acquired a
great importance in the technique of low temperatures and of liquefying
gases. Mostly, it is not used as the sole agency of cooling but in com
bination with adiabatic expansion. Air, oxygen, and nitrogen can be
subjected to the process at room temperature since their inversion
points lie higher. For hydrogen, T> lies above the boiling points of
liquid oxygen and nitrogen so that it can be made amenable to the
i Hoxton, Phys, Rev. 13, p. 438, 1919.
IV 28 THE SECOND LAW OF THERMODYNAMICS 73
process by being precooled by liquid air. Again 7\ of helium is above
the boiling point of hydrogen (20?5 K). Whether these gases are pre
cooled in this way (cascade system) or by adiabatic expansion, the
process they are subjected to in technical applications is the integral
JouleThomson effect, i.e. an expansion from a high pressure (p\) to a
considerably lower one (pz) with an attendant temperature drop from
Ti to T2. These changes are still governed by eq. (4.43) or
x(02, r 2 )  x(i, Ti) = 0. (4.55)
We may ask under what conditions the maximum cooling is
attained, (T% = min.), if we start from a given temperature TI. This
question is, in fact, of practical importance, because the temperature
Ti to which the system can be precooled is usually dictated by technical
considerations. Let us suppose, for a moment, that the end pressure
p2 is also given (Ti = const, p2 = const). Then eq. (4.55) is a
relation between only two variables T^ and pi, and the condition for
T2 = min. is obviously dT^/dpi = 0. The total differential of the
equation is
(IlMl;)*^ <
Hence we conclude that, in order to obtain the lowest end tem
perature, the pressure pi must be chosen so as to satisfy the condition
r = 0. (4.57)
The fact that this condition is independent of p2 shows that our
assumption of a fixed end pressure was unnecessary. The choice of pi
as a root of eq. (4.57) assures the maximum of cooling for any end
pressure, provided the difference Ap is not very small. However,
owing to eq. (4.45), this condition is precisely of the same form as that
for the differential inversion point, which we transformed into (4.49)
(with the sign of equality). For a given (initial) temperature, the
optimum in the integral JouleThomson effect occurs at the same pressure
as the inversion in the differential effect. If the substance follows the
law of Van der Waals this pressure is determined by eq. (4.53), or
resolving it with respect to TT:
(4.58)
Exercise 39. Starting from eq. (4.47), prove the relation
(4.59)
74 TEXTBOOK OF THERMODYNAMICS IV 29
Exercise 40. Using partial differentiation (as in exercise 33) derive the relation
. ,
dp/ v \dp/T \dT/ p \dp
With the help of (3.26), (4.45), and (4.27) transform it into

Exercise 41. Calculate the differential inversion points, for the gases and
pressures of Table 7 (excepting air), making use of formula (4.53) and of the data of
Table 2 (on p. 14). Compare them with the measured values.
29. Realization of the absolute scale of temperatures. The abso
lute temperature T was introduced in section 4 as that which would
be measured by means of an ideal thermometer filled with a perfect
gas as thermometric fluid. This temperature acquired a more general
significance when it was found (section 21) that the efficiency of the
reversible Carnot cycle is independent of the nature of the working
system and, being always the same as in the case of a perfect gas,
can be very simply expressed in terms of T. This fact led, in turn, to
the possibility of defining the concept of entropy by its differential
dS = dQ/T, in which the reciprocal of the absolute temperature
appears as the integrating multiplier.
The role which T plays in the analytical formulation of the second
law makes it a quantity of general thermodynamical importance and
gives a practical way of indirectly determining the absolute tempera
ture, although the perfect thermometer is only an abstraction. In
fact, any thermodynamical equation derived from the second law
represents a relation between T and the other variables of the system
to which it applies. Theoretically it is, therefore, possible to express
the absolute temperature T in terms of measurable characteristics of
any system of nature. On the other hand, one can read directly,
under the same conditions, the arbitrary temperature / on a real
thermometer and, comparing it with T, graduate this thermometer in
the absolute scale. This procedure was conceived and carried out by
Lord Kelvin who used, for this purpose, the equation of the Joule
Thomson process. Let us take the form (4.48) which refers to the
case of small pressure differences Ap. Suppose that we have estab
lished a correspondence between the scales of T and /, so that we can
regard them as functions of each other. We wish to introduce into the
equation the quantities A<, (9v/3/)pi C'P measured by means of the
real thermometer. We obtain immediately, AT" = At*dT/dt,
(dv/dT) p (9v/3/) p dt/dT; moreover, from the definition of molal
IV29 THE SECOND LAW OF THERMODYNAMICS 75
heats we have, in general, c Dq/dT = (Dq/dt)dt/dT = c'dt/dT and,
in particular, c p = c' p dt/dT. The substitution into (4.48) gives
Ap dt
whence
dT __ Qp/30
T
The right side of this expression contains only readily measurable
quantities, so that the absolute temperature can be found by integrat
ing it. In practice it is convenient to measure the temperature t with
a gas thermometer filled with the same gas which is used in the Joule
Thomson process. We have then, according to (1.08), v = i>o(l + a/),
and this is the same v as in the denominator of (4.61), since p is there
considered as constant. Further, (dv/Qf) p = VQOI, so that the integral
of the right side becomes
adt
f
Suppose that to the point / = C corresponds T = TO, then we
can adjust the scales so (compare section 4) that to t = 100 C corre
sponds T = To + 100. The integral of (4.61). is, therefore,
log (r/7o) = K(t), (4.63)
and for t = 100 one obtains an expression for To in terms of measur
able quantities
To = 100[exp X(100)  1]. (4.64)
The number To indicates how many degrees of the absolute scale
the zero point of absolute temperature (T = 0) lies below the point
of freezing water (/ = 0). It is remarkable that the measurements of
A//A/> need not be carried out at very low temperatures, in order to
obtain it, but only between / = and t = 100. In this range the
numerical value of the term c' p &t/vo&p is pretty small. In the case of
helium, it is smaller than 0.001 in any part of the interval, so that
neglecting it altogether would already give an approximation of better
than 0.1%. If this is done, eq. (4.62) reduces to K(f) = log (1 + a/),
whence To = I/a, as if helium were a perfect gas. With the experi
mental a = 0.003659 this gives To = 273.32, which is, in fact, pretty
close to the true value. The most accurate determinations of TO were
carried out by Roebuck (by this method *) and by Henning
1 J. R. Roebuck, Proc. Am. Soc. Arts, Sci. 60, p. 537, 1925.
76 TEXTBOOK OF THERMODYNAMICS IV 30
and Heuse (by a different one *). The weighted mean of their
results 2 is
To = (273.18 0.03) K,
where the symbol K indicates degrees in the absolute or Kelvin
scale. Nevertheless, we shall have to adopt here the old value of
To = 273M on which all current tables are based. A change in the
definition of the zero point has farreaching consequences and can be
made only by international agreement.
As stated in the beginning of this section, any phenomenon which
satisfies the following two conditions can be used for establishing the
absolute temperature scale: it must be (1) expressible by a theoretical
formula as a function of T, and (2) measurable in terms of a relative
scale /. Temperatures in the vicinity of K are being established by
means of the magnetocaloric effect (compare section 136), the same
effect by means of which these extremely low temperatures are reached.
Exercise 42. In the range from to 100 C (at 1 atm) the Van der Waals
equation gives a sufficient approximation for a rough estimate of the third term in
the denominator of the integral (4.62). Calculate its maximum value in this interval
from eq. (4.49) and the data of Table 2, for H 2 and N 2 . Estimate the percentual
error in To, if this term is neglected.
Exercise 43. How can the formula O/a)r = T(dp/dT) v  p be used to
establish the absolute scale of temperatures? Suppose that (d/dt>)*, (dp/dt) v
and p can be measured as functions of the relative temperature t.
30. Statistical interpretation of the entropy concept. Matter con
sists of discrete particles (atoms, molecules, etc.), and the actions
exercised by material systems upon one another can be traced to the
effects of these ultimate parts. Thus the pressure of a gas upon a wall
is due to its bombardment by the moving molecules of the gas which
hit it at very short, but irregular, intervals and transfer to it varying
amounts of momentum. From this point of view, the pressure is a
statistical average which is automatically taken and presented to us by
the integrating action of our measuring instrument, the manometer.
The same is true with respect to other thermodynamical variables and
functions. The temperature, in particular, can be brought in con
nection with the average of the kinetic energy which is distributed
over the individual particles according to the laws of chance or the
laws of probability.
The mathematical theory of probabilities had its origin in the
analysis of games of chance. Suppose that a deck containing an
* F. Henning and W. Heuse, Zs. Physik 5, p. 285, 1921 ; W. Heuse, 37, p. 157, 1926.
* R. T. Birge, Phys. Rev. Suppl. 1, p. 18, 1929.
IV 30 THE SECOND LAW OF THERMODYNAMICS 77
equal number of black and red cards is manufactured and shuffled so
perfectly that the chance of drawing from it a black or a red card is
exactly equal. In other words, if the drawing is repeated a large
number of times Z, the card being returned and the deck reshuffled
every time, the number of red and black cards drawn will be, relatively,
the closer to Z/2 the larger Z is. We say then that the probability of
drawing a red card is one half: PR = ^, and, in the same way, the
probability of drawing a black card PB = ^ What is now the prob
ability PRB of obtaining, in two successive drawings, first a red and
then a black card? There are four possible cases for the result of this
double drawing: (1) red, red; (2) red, black; (3) black, red; (4) black,
black. Since the chance of each of the four cases is the same, we
conclude that the probability P RB = J or P RB = P/ePz*. This is
an example of the multiplicative law of probabilities. In general, the
probability Pi2 of a complex event, which consists of two simple
events, is equal to the product of their respective probabilities PI, P%
Pi2 = PiP 2 . (4.65)
The multiplicative law holds only when the simple events are
statistically independent, meaning by this that the outcome of the first
in no way influences the chances of the second.
The application of the laws of probability to the molecular structure
of matter is treated in the branches of physics called statistical
mechanics and kinetic theory. Here we shall give only a brief refer
ence to some notions about the entropy which are, in part, postulated,
in part, deduced in these disciplines. A striking property of the entropy
is its onesidedness. In any spontaneous process of nature it can only
increase, and the state of equilibrium is the state of maximum entropy
(compare also next section). On the other hand, when we regard a
physical system as an assembly of interacting atoms and molecules,
we say that it will change in the direction of a more probable state, as
to positions and energy distribution of its ultimate units: The final
state will be that of the highest probability. The question arises,
therefore, whether it is possible to give a definition of the probability
of a state which would bring out its parallelism with the entropy and
permit to follow it through in a mathematical way. This is, pre
cisely, what is accomplished in statistical mechanics: The probability
of a state is defined there as the number of ways in which this state can be
realized. Let us illustrate this by the example of pure and mixed
crystals. Consider a crystal consisting of Zi identical atoms in a reg
ular arrangement. We shall discuss here the probabilities, only in so
78 TEXTBOOK OF THERMODYNAMICS IV 30
far as they depend on the normal position of the atoms, and leave out
of consideration the deviations and velocities due to their thermo
kinetic motions. If we could label and individually distinguish the Z\
atoms, every permutation of them would lead to a new way of realizing
the crystal, and the total number of ways would be equal to the
number of possible permutations P s i = Z\ !. This is called by Gibbs
the specific definition of the probability. In modern statistical mechanics
it is recognized, however, that it is not possible to tell the atoms apart
so that all the permutations must be considered as one single realization
Ppi = 1: the generic probability definition which was preferred by
Gibbs himself. On the other hand, let us take the case of a mixed
crystal, that is Z\ atoms of one kind and Zi of another arranged in a
joint lattice. If we assume that every distribution of atoms of the
second kind between those of the first represents still the same macro
scopic state of the mixed crystal, we find that the number of possible
realizations is, in the specific definition, again equal to the permutation
number P s i 2 = (Z\ + Z%)\. In the generic definition it is smaller
because the permutations of atoms of each kind among themselves
(Z\\ and Z2 ! .) do not count. The previous number, therefore, must
be divided by the product Z\ \ Z* \
Let us consider the process of combining two pure crystals (Z\
and Z2) into a mixed crystal. The initial probability of the system,
according to the multiplicative law (4.65), is P\Pi\ We see that it is
increased in the process in the ratio given by the expression (4.66), no
matter which of the two definitions we use. We repeat that the pre
ceding discussion implies that all the realizations which we counted
belong to the same macroscopic state of the crystal. To what extent
this is the case in reality, will be considered in sections 98 and 99.
What is now the mathematical relation between the probability
and the entropy function? The probability of complex systems follows
the multiplicative law, the entropy the additive. We cannot, there
fore, make them simply proportional. But if we take the logarithm
of the probability, the difficulty is removed, since eq. (4.65) then
becomes log Pi2 = log PI + log P%. In view of this, Boltzmann
postulated
S = * log P, (4.67)
as a general principle which now bears his name. Here k = R/n A
is the Boltzmann constant as defined by eq. (1.14). Applying this to
IV 30 THE SECOND LAW OP THERMODYNAMICS 7$
the above example, we find that the entropy increase in the process
of producing a mixed crystal out of two pure ones is
AS = k [log (Zi + Z 2 )!  log Zi!  log Z 2 !]. (4.68)
As stated above, the multiplicative law of complex probabilities
applies only when the simple events are statistically independent.
On the other hand, the entropy of a complex system is additive, only,
when the equilibrium of its constituent parts is characterized by a
uniform temperature (section 22). It follows from this that the
thermodynamical criterion of two statistical systems not being inde
pendent is that they come in equilibrium with different temperatures
(compare footnote on p. 60).
This is not the place to enter into the applications of Boltzmann's
principle. We shall only mention one or two of its immediate conse
quences which will be useful in interpreting and rounding out the
thermodynamical treatment. We restrict ourselves to the case when
the events are considered as statistically independent.
(A) There is an alternative formulation of Boltzmann's principle
which we give only for two special cases as we shall not need it in its
generality. (1) If the particles of a system move in a field of conserva
tive forces, the probability of finding a particle in the space element
Ar is
Pdr = Cexp (e po t/&7>r, (4.69)
where e po t is the potential energy of the particle, when in the space
element dr, and C is independent of the coordinates of the particles.
(2) A similar relation exists for the total energy, its formulation is
particularly simple in the quantum theory. If a system consists of
identical elements (atoms, molecules, linear oscillators, etc.), each
capable of assuming the quantum states 0, 1, 2, ... with the respective
energies eo, ei, 62, . . ., the probability of finding an element in the
state (j) is
P,= C'exp(e,/tr). (4.70)
(B) In classical statistics (when quantum restrictions do not exist
and all energy levels are permissible) Boltzmann's principle leads to
the equipartition of energy. The mean kinetic energy of an element
is equal to kT/2 per degree of freedom:
Sun = nkT/2, (4.71)
n being the number of (translational and rotational) degrees of
freedom. As to the potential energy, this law applies to it only in
80 TEXTBOOK OF THERMODYNAMICS IV 30
the case when the forces producing it are elastic forces (i.e. are pro
portional to the distance).
Exercise 44. A coin is tossed three times; what are the probabilities of (1) head
being up all three times, (2) head being up twice and tail once?
Exercise 45. Calculate from the formula (4.69) the relative number of mole
cules of air in a space element dr of the earth atmosphere at the height h above the
surface of the sea. Use a mean molecular mass (m), and, substituting k R/n At
show that one obtains in this way the barometric formula of Laplace.
CHAPTER V
THE TREND OF THERMODYNAMICAL PROCESSES
TOWARDS EQUILIBRIUM
31. The direction of spontaneous processes. Let us consider a
complex thermodynamical system whose parts may or may not be in
equilibrium. We assume it prevented from heat exchange with the
outer world by an adiabatic envelope but capable of doing work. If
all its microscopic dynamical, electrical, etc., parameters were given,
as well as the outer forces acting on it, its future states would be deter
mined and predictable. Can we foresee anything about the spon
taneous processes in it and their direction, knowing only its thermo
dynamical characteristics? The two laws of thermodynamics as
stated in the preceding chapters will not lead us very far in this
endeavor : the first law only states that one type of energy is converti
ble into another, but, as far as this principle is concerned, the con
version can go either way, provided the geometrical constraints of the
system permit it. The second law tells us a little more about the
direction of spontaneous processes: only such changes are permissible
in which the entropy does not decrease. This excludes, at once, a
large number of conceivable occurrences as incompatible with this
law. Still, the second law is incomplete in that it leaves the question
open whether the entropy actually will increase when conditions per
mit it. In fact, in the enunciations of the preceding chapter the fun
damental fact is not contained that there is in every thermodynamical
system a tendency towards a definite state of equilibrium.
Fortunately, the empirical evidence relating to these questions can
be summarized in the form of two simple rules supplementing the first
and the second law of thermodynamics. In the first place, the internal
energy U has the nature of a potential energy. A dynamical system
which is at rest, at a certain moment, will set itself in motion if its
constraints permit a decrease of the potential energy. The potential
energy has a tendency to diminish and to be converted into kinetic
energy, so that equilibrium can ensue only when it has reached its
minimum. In a similar way, the internal energy of an isolated thermo
dynamical system tends to decrease when the existing conditions give
81
82 TEXTBOOK OF THERMODYNAMICS V 31
it a chance of being converted into forms of energy of nonpotential
nature. U tends towards its minimum, and the system will not be
in equilibrium until this minimum is reached.
U>U min . (5.01)
In the second place, the entropy of isolated systems tends to increase.
Processes attended by an augmentation of the entropy are not only
permissible, as the second law states, but one of them will necessarily
take place spontaneously. Equilibrium will be reached only when S
has attained its maximum
SS ma *. (5.02)
This condition can be regarded as a portion of the second law, a
third part of the entropy principle, of equal importance with the first
two enunciated in section 23. We speak of it separately in this place,
in order to bring out its analogy with the rule (5.01) supplementing
the first law.
From this discussion follow the necessary and sufficient conditions
of equilibrium of a thermodynamical system : The internal energy must
have its minimum and the entropy its maximum value. It is known,
from mathematics, that this will be the case if, for all virtual infinitesi
mal changes of state, consistent with the nature of the system, the
variations 6 U and SS vanish :
5 [7 = 0, 55 = 0. (5.03)
There is a difference between the differential and the variation.
The differential dU refers to an actual infinitesimal change of state
which the system undergoes in a specified real process. In thermo
dynamics, the initial and the infinitely close final state must be states
of equilibrium if they are to define a differential. On the other hand,
the variation has reference to all conceivable (" virtual' ') changes
whether they conserve the equilibrium or not. A variation may start
from a state of equilibrium and lead away from it. It may even con
nect two (infinitely close) thermodynamically unstable states. The
reason for such a definition of the variation is that we do not know in
advance which virtual changes are permissible, and we wish to select,
a posteriori, those which are compatible with the conditions (5.03)
and with other conditions, inherent in the nature of the system and in
the concept of equilibrium. It is clear from this explanation that the
virtual changes are the more comprehensive class: every actual change
is contained among the virtual changes, but the reverse is not true.
V32 THE TREND TOWARDS EQUILIBRIUM 83
32. Variation of the internal energy. Partial molal quantities.
The systems and states we are going to deal with must be capable of a
thermodynamical description. We have to restrict our considerations
to systems which can be divided into phases with the following prop
erties: though, possibly, not in physical or chemical equilibrium, each
phase must be uniform with respect to temperature, pressure, and
composition so that it has a definite internal energy and entropy. For
all virtual changes in which no matter is added to the system, holds
the fundamental formula (4.07), or
dU = T5S  pdV. (5.04)
We can choose the volume V and the entropy S as the thermo
dynamical variables. In addition, the system may be characterized
by other parameters specifying its mass, chemical composition, etc.
To fix our ideas let us assume that the system is not chemically pure
but a mixture of Ni mols of a first chemically pure component, N% mols
of a second, and so on. The internal energy is then a function of the
variables
The variation of U is the most general change which takes place
while all the variables undergo virtual changes,
We shall only consider virtual changes in which no matter is added
to the system so that the variations dN k represent increases of some
components at the expense of other components. In this case the
formulas (5.04) and (5.06) are valid simultaneously and apply to all
possible virtual changes satisfying this condition. In particular, we
may consider the case when the composition remains altogether
unchanged, SN k = (compare, in this connection, section 39). Com
paring the two equations, we find
(H)  T  (If) > < 5 ">
\OO/V.N \OV/S.N
The subscript N is usually omitted because the symbol of partial
differentiation implies that all the other variables are to be kept con
stant. Substituting this into the general eq. (5.06) and subtracting
(5.04), we obtain the remarkable relation
84 TEXTBOOK OF THERMODYNAMICS V32
These results can be generalized for systems with several phases
(provided T and p are uniform throughout), but we need not deal with
this aspect of the question here since it will occupy us in section 42,
in connection with the problem of equilibrium. The partials of
eq. (5.08) should not be confused with the partial molal internal ener
gies which are defined as follows :
dNk/^r
In the case of a chemically pure substance, U = Nu, where the
ordinary molal internal energy u is a function of temperature and pres
sure only and remains constant when p and T are kept constant.
Therefore, Uk is a generalization of u for mixtures and solutions,
becoming identical with it in the extreme case of pure substances.
Another partial molal quantity worth mentioning is the partial molal
volume. When several substances are mixed or dissolved in one
another, the total volume is, generally, not additive but a complicated
function of the mol numbers (compare also section 57)
V = V(p, T,NL.. N 9 ).
The partial molal volume of each component is defined as
/oT7\
(5.10)
The relation of the functions (3 U/dN k ) s . v to partial molal quan
tities will become clear in the next sections.
Exercise 46. The measurements by Wade (J. Chem. Soc. 75, p. 254, 1899) on
the volume of aqueous solutions of NaCl can be represented by the following empiri
cal formula (due to Lewis and Randall)
where N\ refers to water and Nt to NaCl. Calculate the volumes of solutions with
0.1, 0.3, 0.5 mol of NaCl per 1000 g of water and the partial molal volumes of NaCl
in all these cases.
Exercise 47. Prove for the perfect gas the formula
Directions. Calculate &U = d(Nu) from (3.18) and eliminate ST and Sv with
the help of the conditions 55 5(Ns) = and dV  5(Nv)  0. For s, use the
formula (4.17).
V33 THE TREND TOWARDS EQUILIBRIUM 85
33. Construction of new characteristic functions. Further prop
erties of the heat function. In addition to the characteristic functions
U and 5, we had occasion to introduce in section 17 the heat function
defined as
X  U + pV, (5.11)
while its differential, in the case of a simple system, is obtained from
(3.22) and (4.07),
dX = TdS + Vdp, (5.12)
whence in an isobaric process (dp = 0)
Q P = X 2  Xi. (5.13)
Instead of using the variables 5, V of the preceding section, we
can describe the system also by the variables S, p, and use X as the
appropriate characteristic function. The term pV in (5.11) has the
effect that its differential d(pV) = pdV + Vdp cancels the term
pdV of (5.04) and leaves in its place Vdp. In this way the differ
ential dp is introduced instead of dV. In mathematics the method of
replacing a term ydx by the term xdy by means of subtracting the
differential d(xy) is called a Legendre transformation. It is obvious
that we can construct by this expedient still two more characteristic
functions from the expression (5.04), namely, subjecting to Legendre
transformations the first term, or both terms simultaneously. We
shall discuss these functions in the next sections and reserve the
remainder of this one for a few remarks about the variation of the heat
function.
As explained in the preceding section, we assume that the state of
the system depends on the mol numbers Ni. NZ, . . . N ff in addition to
the thermodynamical variables:
X = X(5, p, Ni,... N ff ). (5.14)
We can obtain, for the variation of X, two expressions closely
analogous to those for dU in the preceding section. On the one hand,
the relation (5.12) is valid also for the variations
dX = TdS + Vdp.
On the other hand, we find from (5.14)
86 TEXTBOOK OF THERMODYNAMICS V33
As in the preceding section, we conclude by comparing the two
expressions
(i)  r  (HD  v 
\dS/p,N \3P/S.N
(The subscript N is usually omitted). Another form of the variation
of (5.11) is dX  Vdp = dU + pdV. Hence and from (5.06) and
(5.15)
We cannot conclude from this equation that the terms are equal
individually:
9 X\
because the virtual changes 6N k are not independent: we have excluded
changes in which matter is added to the system. However, the proof
of the identity (5.17) can be easily supplied by partial differentiation.
In fact, let us describe U by the coordinates S, p instead of 5, V. We
can effect this change of variables by substituting for V its expression
in terms of 5, p, N. We need not bring S in evidence since it is con
sidered as constant, and we write instead of (5.05)
U  U[V(p, N), N].
By the rules of partial differentiation
According to (5.07), the second term on the right is [d(pV)/dN] St p .
Taking it to the left side, the equation becomes identical with (5.17).
Exercise 48. Consider the general form of d U based on (2.02)
Show that the number of characteristic functions which can be constructed from it
by Legendre transformations is 2 n + l 1.
Exercise 49. Prove that, for a perfect gas,
Directions. Calculate dX = 8(N\) from (3.28) and eliminate ST with the help
ol the conditions 6S d(Ns) and dp = 0. For s, use the formula (4.18).
V34 THE TREND TOWARDS EQUILIBRIUM 87
34, The work function. We subject to a Legendre transformation
the first term of the differential of the internal energy (4.07)
dU = TdS  DW. (5.18)
Subtracting from it d(TS), we find
D*   SdT  DW, (5.19)
where
*  U  TS. (5.20)
The function ^ is useful in bringing out certain properties of iso
thermal processes. We have, in this case, dT = and
d& =  DW. (5.21)
Integrating this over a reversible process at constant temperature
from the initial state (1) to the final (2),
*!  * 2 =  A* = W. (5.22)
The work done by a system in a reversible isothermal process is equal
to the decrease of the function V. The case when the system remains
in contact with a heat reservoir of constant temperature but the
process is not conducted reversibly is sometimes, loosely, called an
irreversible isothermal process. In this kind of a process the tempera
ture of the system is not strictly equal to that of the reservoir and is,
in fact, not strictly uniform. We have seen in section 12 that the
process is less efficient and the work done in it by the system is less
than in the reversible case. We can, therefore, say that ^i 2
represents the maximum work that can be obtained from the system in
any isothermal process leading from the state (1) to the state (2).
Applying eq. (5.20) to the initial and final states and taking the
difference,
Ui  E/2  *i  * 2 + T(Si  S a ). (5.23)
The second term on the right side is, according to eq. (4.11), equal
to Q, representing the heat transferred to the heat reservoir during
the process. Helmholtz, who used the function ^ a great deal, called
it the free energy because the difference ^f\ ^2 represents that part
of the decrease of the internal energy U\ Uz in the process which
is available for work while the bound energy T(S\ 2) is lost to the
reservoir maintaining the constancy of temperature. However, the
name cannot be regarded as very fortunate because it loses sight of
the fact that there exist in nature endothermic processes in which
T(S\ ^2) is negative. In this case, work is done by the system, in
88 TEXTBOOK OF THERMODYNAMICS V 34
part at the expense of the heat reservoir, which supplies heat instead
of receiving it. The decrease &i ^2 can, therefore, be larger than
Ui  C/ 2 , as was first pointed out by Gibbs. To make matters worse,
in a part of the chemical literature of this country the name "free
energy" is applied to a different function (namely, to $ defined in the
next section). We think it, therefore, advisable to avoid the term
"free energy" altogether and shall refer in this book to the function ^
as the work function. This name was introduced by R. H. Fowler 1
and is highly appropriate in view of the physical meaning of the
quantity ^ as expressed in eq. (5.21). This relation means that, as
long as the temperature is kept constant, the elements of work used in
thermodynamics are exact differentials which possess the potential &.
We shall discuss the very important question, under what circum
stances ^ tends to decrease, in section 36 and shall say here only a few
words about its partial derivatives. In the case of a simple system,
DW = pdV, and the variation becomes, in accordance with (5.19),
6* = SdT  pdV. (5.24)
The appropriate coordinates for the description of the work func
tion are, therefore, T and V in addition to the mol numbers N k , deter
mining the composition of the system:
V, N l9 ... NJ.
In analogy with the preceding sections, we find
'),
(<
\dTJv, N \OV/T,N
& P.
(The subscript N is usually omitted). As to the partials with
respect to Nk, it is easy to prove the identity (exercise 51)
y
by means of which eq. (5.08) can be expressed in terms of derivatives
of the function ^.
It will be useful to give here the explicit expressions of the molal
work function $ for perfect and Van der Waals gases. We take the
more general case of the formulas (3.19) and (4.20) when the molal
heat is not considered as constant. In the case of the perfect gas,
* '(D  RTlogv, (5.27)
1 R, H. Fowler, Statistical Mechanics, p. 96. 1929.
V35 THE TREND TOWARDS EQUILIBRIUM 89
of the Van der Waals gas, according to (4.26) and (4.28),
* = co'(r)    RTlog(v ft), (5.28)
where o/(r) is an abbreviation for
'(r) y cjT  rj c ^ + */<> TS' Q . (5.29)
Exercise 50. Prove, for the perfect and for the Van der Waals gas, the relation
Directions. Calculate 5* = d(Ntf with ST = from (5.27) and (5.28). Elimi
nate 5v by means of the condition 5 V = d(Nv) = 0.
Exercise 51. Prove the identity (5.26) in the following way: In (5.05) replace
5 by a function of V, T, N, i.e. U  U[S(T, N), N], and take the partial (dU/dN) T . v
in analogy with the procedure at the end of section 33.
Exercise 52. Consider the case of a system with many degrees of freedom
when the element of work has the general form (2.02): DW = 2y k dX k . Write out
the differential of the work function * = U  TS and, by applying the reciprocity
relations (2.10), prove the equation
i) . (5.30)
Jx
35. The thermodynamic potential. Let us consider a system
which depends on some nonmechanical (e.g. electrical, magnetical,
etc.) variables, in addition to p and T. The element of work done by
it can, then, be represented as DW = pdV + DW ', where DW is the
work done by the nonmechanical forces. Accordingly, the differen
tial (5.04) of the internal energy takes the form
dU = TdS  pdV  DW. (5.31)
We subject the first and second term on the right side to Legendre
transformations adding  d(TS) + d(pV). The result is
d$ =  SdT + Vdp  DW, (5.32)
where
$ = U  TS + pV (5.33)
is known as the thermodynamic potential,
In the laboratory, processes (like chemical reactions, etc.) are very
often conducted at constant temperature and constant pressure.
The temperature is maintained by a suitable heat bath while the pres
$0 TEXTBOOK OF THERMODYNAMICS V 35
sure is that of the terrestrial atmosphere. Under these conditions
(dT = 0, dp = 0), eq. (5.32) reduces to
' d*   DW, (5.34)
or integrating this over a reversible path (at T =* const, p = const)
from the initial state (1) to the final (2)
$1  $2 =  A$  W. (5.35)
>
The nonmechanical work done by a system in a reversible isothermal,
isobaric process is equal to the decrease of the function <. When the
process is led between the same two states in an irreversible way
(i.e. when the temperature and pressure are imperfectly maintained),
it is more wasteful. The difference $1 $2 gives, therefore, the maxi
mum of the nonmechanical work that can be obtained from the system in
any process leading from state (1) to state (2) of the same temperature
and pressure.
The great importance of the thermodynamic potential arises from
the fact (which we shall prove in the next chapter) that, in the equilib
rium state of a heterogeneous system, p and, T are the same in all its
phases. They are, therefore, the most convenient thermodynamical
variables, and eq. (5.32) shows that & is the appropriate characteristic
function for this choice of variables. In the special case of a simple
system DW f = 0, and the variation becomes
*$ =  S5T + V6p, (5.36)
while $ has the expression
In analogy with the preceding sections
(5.37)
dp/T.n
(The subscript N is usually omitted). The partial derivative with
respect to Nk must be taken at constant p and T: it is, therefore, the
partial molal thermodynamic potential, according to the definition of
the partial molal quantities given in section 32.
(538)
V36 THE TREND TOWARDS EQUILIBRIUM 91
It is easy to prove the identity (exercise 54)
or from (5.17) and (5.26)
We shall have occasion to use the expression for the molal thermo
dynamic potential of a perfect gas. From (5.33), (4.20), and (3.19)
(5.41)
where
(5.42)
In the particular case c p = const,
co(r) = c p T(l  log r) + uo  TSQ. (5.43)
Exercise 53. Derive the explicit expression of <f> for the Van der Waals gas:
V"(r)+*r log (/>+") +^~ (5.44)
\ V l / V V
Exercise 54. Prove the identity (5.39), making in exactly the same substitu
tion for V which was made at the end of section 33 in U (to prove 5.17).
Exercise 55. Consider the case of a system with many degrees of freedom when
the element of work has the general form (2.02): DW HyidXk. Write out the
differential of the generalized thermodynamic potential & = U TS + Sy^Yjfc, and
prove, by applying the reciprocity relations (2.10), the equation
36. The decreasing tendency of the thermodynamic potential and
of the work function. We have found in eqs. (5.22) and (5.35) that
work is done by a system (at T = const) at the expense of the work
function, and nonmechanical work (at T = const, p = const) at the
expense of the thermodynamic potential. The question arises, how
ever, as to the tendency of these functions to decrease: if the differ
ences 1 ^2 or $1 $2 are available in a system, will they be spon
taneously converted into work, and under what conditions?
Let us start our considerations from the work function ^ U TS.
It is obvious that ^ would decrease (if possible) in a system " left to
itself " or " isolated", but remaining at constant temperature. In fact,
92 TEXTBOOK OF THERMODYNAMICS V36
we know from section 31 that the positive term U has a tendency to
decrease and the negative TS to increase (in absolute value). How
ever, in the processes which we wish to consider the systems are not
isolated: in the first place, there is the heat bath by means of which
the temperature is maintained; in the second, the environment may
do work against it through the forces of pressure. Let us try to make
the system isolated by including in it both these influences. With
respect to the heat bath, this is easy since it does not do any work of
either a mechanical or a nonmechanical nature. According to eqs.
(5.21) and (5.34), the values of the functions < and ^ characterizing
the heat reservoir remain constant in view of this. We may regard it,
therefore, as a part of the system without changing anything in the
available differences ^i ^2 and 3>i <i>2, or, in other words, we
may base our discussion on ^ = U TS of the enlarged system (with
inclusion of the heat bath). The conditions are simplest when other
interactions with the outer world are precluded. This is the case when
the volume of the system is constant, for instance, when it is surrounded
by a rigid heatconducting envelope and immersed in the heat bath:
when the temperature and the volume of a system are kept constant, its
work function has a tendency to decrease. Any spontaneous isothermal
and isochoric process goes on until ^ reaches its possible minimum.
The necessary and sufficient condition for the starting and proceeding
of a process is that the work function must be larger in the initial
state than in the final (^i ^2 > 0). This can be used as a criterion
whether a chemical reaction will go. At constant temperature
and volume the maximum possible change of V is a measure of the
chemical affinity of the substances entering into the reaction. It is to
be noted that the work done by this reaction, if any, is nonmechanical
work: from dV = 0, there follows DW = DW.
The conditions are different when the pressure is maintained con
stant, instead of the volume. The practical method of accomplishing
this is to subject the system to the hydrostatic pressure either of the
atmospheric air, or of another medium (liquid or gas) compressed by
a suitable load. If the system expands under this pressure from the
initial volume V\ to the final V^ the work of expansion p(V^ Vi) is
converted into the potential energy of lifting the load or the air column
resting on the system, and pV is a measure of this potential energy at
any moment. In addition to 7, having the nature of a potential
energy, we have here still the item pV which tends to be converted
into energy of a nonpotential character. We have, therefore, to use
the thermodynamic potential $ = U TS + pV: including the heat
bath into the system as in the former case, we see that this function has
V37 THE TREND TOWARDS EQUILIBRIUM 93
a tendency to decrease (at T = const and p = const), since the two
positive terms tend to decrease, the negative to increase (in absolute
value). The criterion for a process taking place under these conditions
is that the thermodynamic potential must be larger in the initial state
than in the final ($1 $2 > 0). This applies in particular to chemical
reactions: they will continue as long as they can produce nonmechani
cal work. The measure of chemical affinity in a reaction conducted
in an isothermal and isobaric way is the maximum nonmechanical
work W = $1 $2 which can be obtained from it.
All three terms of $ tend to decrease, but it would be wrong to
conclude that they always diminish separately. The functions U, S, V
are all interdependent, and the reduction of one may force the growth
of another. Only the sum <i> always decreases, each individual term is
only then sure not to increase when the other two remain constant.
From (5.35) and (5.33)
W' =  A* = (Ui  U 2 )  r(Si  5 2 ) + p(Vi  F 2 ). (5.46)
The term T(S\ 62), representing the heat imparted to the heat
bath, is often negative (endotermic reaction), and the third term
p(Vi 2) is also sometimes negative (reaction with expansion).
In later chapters we shall have to consider the case when the sys
tem is in partial equilibrium, its phases having the same temperatures
but standing under different pressures pi, p2, p a * Since U and 5
are additive, the thermodynamic potential has, then, the form
* = U  TS + piVi + . . . + p a V a . (5.47)
It is obvious that it tends to decrease also in this case, because the
terms pjVj are the potential energies due to the expansion of the indi
vidual phases.
37. The GibbsHelmholtz equation. Galvanic cells. Another
form of the relation (5.46) is
A$ = X 2  Xi  T(S 2  Si),
or substituting for S from eq. (5.37)
This relation is known as the GibbsHelmholtz equation. It applies
to any isothermal difference A< as it is based only on the definitions of
the functions $ and X. In the special case of an isothermalisobaric
94 TEXTBOOK OF THERMODYNAMICS V37
process, we have A$ = W, and from (4.36"), AX = Q p (isobaric heat
of reaction),
(s  49)
The left side of the equation is identical with T(S2  Si) and
represents the energy gained by the system from the heat reservoir,
in the isothermalisobaric process we are considering. It is interesting
to note that this heat is proportional to the temperature coefficient
of W. The GibbsHelmholtz equation has many uses; as an example
we mention here its application to the electromotive forces of galvanic
cells. The active substance in a galvanic cell is an electrolyte whose
negative ions react with the material of the positive plate, transferring
their electric charges to it. At the same time, an equal number of
positive ions is deposited upon the negative plate and neutralized by
the (negative) current flowing to it through the outer circuit of the
cell from the positive plate. According to Faraday's law, the electric
charge supplied to the circuit by 1 mol of the electrolyte is <rF, where
F is Faraday's equivalent, (9648.9 0.7) abs em units, and o the
valency of the ion. The electric energy w f of the current (per 1 mol
of the electrolyte) is obtained multiplying this by the electromotive
force E,
w r = 0FE, (5.50)
or substituting into (5.49)
r (af)  E + >/** < 5  51 >
where q p is the heat of reaction, referred to 1 mol. When q p is
expressed in calories and E in international volts, the numerical value
of the Faraday becomes F = 23 046.
Let us apply this to the Weston cell in which the electrodes consist
of Cdamalgam and of pure mercury. It contains the following sub
stances:
HgCd  CdS0 4 (solid)  CdS0 4 (sat. sol.)  Hg 2 S0 4 (paste)  Hg
arranged as shown in Fig. 11. The following empirical formula has
been found (and internationally adopted) for the electromotive force
of the Weston cell
i [1.01827  4.06(/  20) X 10" 5  9.5(*  20) 2 X 10~ 7
+ (20  O 3 X 10~ 8 ] int volt. (5.52)
V37
THE TREND TOWARDS EQUILIBRIUM
95
Saturated Solution
of Cadmium Sulphate
Mercurous
Sulphate
Cadmium
Amalgam
Platinum Wire
Equation (5.51) permits to calculate the heat of the reaction
(Cd + Hg 2 SO 4 = 2Hg + CdSO 4 ) in the cell. It gives, with <r  2,
q p = 47 482 cal mol"" 1 , while direct thermochemical measurements
yielded 47 437 cal mol" 1 . In the Weston cell the temperature coeffi
cient of E is negative and represents an energy loss. There are, how
ever, other cells (see exercise 56) where it is positive, so that a part of
their energy and electromotive force is supplied by the heat bath
(surrounding air).
It is clear from this discussion that in a galvanic cell the heat Q p
of an exothermic reaction is converted into (electrical) work in a
practically complete way. Its
operation is not cyclic but can
be made continuous, [since the
working substances can be fed
in, at the positive plate, and
the waste products taken out,
at the negative. If the same
reaction were used for heating
the boiler of a cyclic heat engine
(section 21), not more than the
fraction (7\  T 2 )/Ti of the
heat Q p could be utilized, ac
cording to eq. (4.03). In view
of this superiority, there was a
great deal of speculation whether
it is feasible to construct a galvanic cell working with the most com
mon industrial fuel, coal (i.e. carbon). But it proved impossible
to bring carbon into electrolytic solution.
For some of the applications it is useful to integrate eq. (5.49) and
to express W = A< in terms of Q p . An integrating multiplier
of the equation is 1/r 2 , and the multiplication by it gives
[d(W'/T)/dT] P = &/r 2 ,
whence by partial integration
)dT
y (5.53)
Exercise 56. The PbHg cell is constructed as follows:
Pbamalgam  PbSO 4 (solid)  NajSO 4 (solution)  HgjSO* (solid)  Hg
Its electromotive force is
E  [0.96466 + 1.740  25) X lO" 4 + 3.8(*  25)* X 10~ 7 ] int volt
Platinum Wire
FIG. 11. The Weston galvanic cell.
96 TEXTBOOK OF THERMODYNAMICS V37
What part of E is supplied by the heat reservoir, and what is the heat of reaction at
25 C?
Exercise 57. In the CuHg cell
HgCu  CuSO 4 (solid)  CuSO 4 (sat. sol.)  Hg 2 SO 4 (paste)  Hg
the molal heat of reaction is, at 20 C, q p =  39 596 cal mol" 1 , the electromotive
force E = 0.3500 int volt. Find the temperature coefficient of E.
Exercise 58. In the same way as (5.48) is derived for A4>, derive for * the
equation
.
Hence prove the relation
Y
*  T f^ dT + Tf(P).l (5.55)
Remark. In the differentiation and integration with respect to T the pressure
is regarded as constant. Therefore, the integration constant may depend on p.
In (5.53) the function of p cancels out when the isobaric difference $2 $1 is formed.
Exercise 59. Prove from (5.20) and (5.25) the analogous equations
(5.56)
~\ \J *. r y Jt I (J J. M f
and
dT + TMv). (5.57)
Forming the difference A = 2 1 leads to the relation
T (^f) v =*** u ' (5.58)
which sometimes is also given the name of a GibbsHelmholtz equation.
CHAPTER VI 1
GENERAL CONDITIONS OF EQUILIBRIUM OF
THERMODYNAMICAL SYSTEMS
38. The method of virtual displacements. The theory of thermo
dynamical equilibrium was developed by Gibbs and patterned by him
after the mechanical theory of statics of Lagrange. We have explained
in section 31 what virtual changes are, but it will be well to say here
a few words about how they are used. Let the parameters of the
system be denoted by 1, fe, n , which may include as well the
thermodynamical variables as those defining the composition of the
system; and let the equilibrium be determined by the condition that
the variation of a certain characteristic function F(i, . . . { n ),
vanish for all virtual changes dlk consistent with the conditions of the
system.
If the parameters are not subject to any further restrictions, the
virtual changes are entirely independent of one another and can be
chosen in any arbitrary way. In particular, we could choose all but
one of them equal to zero ($* ^ 0, and when i ^ k, $ = 0). Then
the condition (6.01) is reduced to (3F/3{*)%  Oi or
 = 0, (*  1, 2, . . n). (6.02)
3&
Since 5* was selected at random, this must hold for each, of the
parameters. We have, therefore, n equations from which the equilib
rium values of the n parameters & can be determined, and this consti
tutes a complete solution of the problem.
However, usually, the parameters are not quite independent but
subject to constraints or subsidiary conditions which take the analyti
1 Suggestion to teachers. Chapters V and VI are both somewhat abstract. It is,
therefore, recommended to interpose between them sections 134, 135, and 136 of
Chapter XX, which contain applications of the concepts introduced in Chapter V.
98
TEXTBOOK OF THERMODYNAMICS
VI 38
cal form of equations imposed upon them. Suppose that there exist
m (< ri) subsidiary conditions
/ifti, ...)
/2i, . . 6.)
0,
0,
/mftl, ...*) 0.
(6.03)
In this case, only n w of the f A; are independent and can be chosen
arbitrarily. The remaining m parameters are determined by these
through eqs. (6.03). The same is true of the virtual changes 5 fc : only
n m of them are within our free choice. In order to find the final
conditions of equilibrium, it is necessary to eliminate the m dependent
virtual changes from the variation dF and to represent it as a linear
form of the independent $ *, only. This could be done by eliminating
m of the parameters from the function F, with the help of eqs. (6.03),
before taking its variation. However, this is seldom convenient, and
it is preferable to take 5F in the form (6.01) and to eliminate the
dependent dfa from it, using a method worked out by Lagrange. The
subsidiary conditions can be, also, thrown into a variational form; in
fact, they apply as well to the values of the parameters & as to &
Mti
= 0.
(6.04)
The difference between (6.04) and (6.03) is, for infinitesimal
changes,
v<
f& = 0, (* = 1,2, . . . m). (6.05)
Often the subsidiary conditions are given directly in the variational
form
(6.06)
which cannot always be integrated. However, it is irrelevant for the
further treatment whether the form is (6.05) or (6.06). To fix our
ideas we use eqs. (6.05) and multiply them, respectively, by Lagrangean
multipliers A,(i, . . . () and add them to eq. (6.01)
w i At
VI 39 GENERAL CONDITIONS OF EQUILIBRIUM 99
The m functions \i can be chosen in such a way as to make the m
first parentheses of this equation vanish
f or k = 1,2, ... m. The remaining expression
Z" ( dF + i 9 ^+ LX 9 
\*t + Xl 5T + ' ' ' + Xm ~
*m+l ^3** 3*
contains now only n m independent $*, so that the problem is
reduced to the case of independent virtual changes treated above: this
equation is satisfied only then when all the coefficients in parentheses
vanish. The conditions (6.01) obtain, therefore, for all values of
k( 1, 2, . . . n). There are n such equations, and together with
the m eqs. (6.03), we have m + n relations between the variables
1, . . . n; Xi, . . . X m , i.e. the number necessary to determine all of
them and to obtain a complete solution of the problem of equilibrium.
Exercise 60. A heavy material point moves in the field of gravity of the earth.
It is constrained to remain in the curve (ellipse)
(ax + Py + yz) 2 + ey* = 1, ax  by = 0,
where x t y, z are cartesian coordinates, z being the vertical. The point comes to rest
when its potential energy has its minimum (z min.) What is its position of rest?
(Use Lagrangean multipliers).
Exercise 61. Make the same calculation for a point which is constrained to
move in the surface
x* H (y cos a + z sin a) 2 f 2p(y sin a z cos a) = 0.
39. Auxiliary constraints. Homogeneity of the thermodynamical
functions. It follows from the considerations of the preceding sections
that the equations of equilibrium (6.07) hold for any system of virtual
changes $& satisfying the main condition (6.01) and the subsidiary
(6.05). In some cases, it is useful to treat the problem of equilibrium,
not in its entirety, but only in part, and to select a special system of
variations 5& compatible with those conditions. It is trivial that the
requirements (6.01) and (6.05) are satisfied if all the virtual changes
are taken equal to zero (3& = 0; k = 1, 2, ...). If only some
arbitrarily chosen of them are equal to naught (provided the subsidiary
conditions permit it), the result is quite consistent with the general
solution but represents only a part of it. In fact, in this case, the
100 TEXTBOOK OF THERMODYNAMICS VI 39
corresponding terms of the expressions (6.01) and (6.05) vanish: the
method of the Lagrangean multipliers leads, then, simply to an incom
plete set of eqs. (6.07).
More generally, we can say that it is permissible to impose upon the
parameters any additional constraints which do not disturb the equilibrium
of the system. At first sight it may appear that this principle is of
little practical use because, apart from such simple cases as that just
mentioned, we must first know what the state of equilibrium is before
we can judge whether a constraint is compatible with it or not. In
practice, however, the simple cases are the most important ones, and
frequently the knowledge of the nature of the system and of the
physical circumstances of its equilibrium gives enough criteria to
decide this question. In fact, the introduction of auxiliary constraints
is a very useful and widely applied expedient.
In this chapter we are going to consider a system consisting of
several phases, each described as well by thermodynamic variables as
by the mol numbers NI, N 2 , . . . N, defining its composition (compare
section 32). We shall make use of auxiliary constraints in two con
nections. The first application consists in breaking up the problem
of equilibrium into two partial problems by first considering the case
when the composition remains unchanged and dealing with the ther
modynamic variables alone. Consequently, the auxiliary constraints
take here the simple analytical form dN\ = $N2 . . . = dN ff = 0. It
is obvious that this assumption is compatible with the notion of
equilibrium because the most perfect equilibrium obtains then when
no process at all is going on in the system. The conditions, the
parameters Nk are subject to, are found separately as the second par
tial problem into which the treatment is broken up.
In the other application the auxiliary constraint takes a geometrical
form. Let us imagine that, anywhere, in the system a thin, closed
layer of its substance is removed and replaced by a rigid and adiabatic
envelope. Will the presence of this envelope disturb the equilibrium?
Strictly speaking, it will. The state of a substance, at any point, is
influenced by all the molecules in a small radius around it (radius of
molecular action). Therefore, a very thin layer of matter, immediately
adjacent to the imaginary envelope, will find itself in a new condition,
and this will influence the system as a whole. However, when the
volume within the envelope is sufficiently large the influence of the
surface layer can be neglected. In fact, the volume increases with
the third power of the linear dimensions, the surface with the second,
so that the influence of the surface becomes relatively less and less
important as the size increases. We shall treat the properties of the
VI 39 GENERAL CONDITIONS OF EQUILIBRIUM 101
surface layer separately in Chapter XII, and we are going to assume
here that all the phases of the system are so large that the surface
effects can be neglected (as well at the boundary as at the surfaces of
discontinuity between the phases). When this assumption is made,
the equilibrium is not disturbed by the rigid envelope just mentioned,
whether we think it surrounding the whole of the system or a part
of it. It is inherent in the notion of a system in equilibrium that
there is in it no relative motion of the parts and no heat transfer
between them. The imaginary envelope will not interfere, therefore,
with any actual process going on in the system. As to virtual changes,
some of them are precluded or restricted, but no new virtual changes
are made possible, by the presence of the envelope, which did not exist
without it. According to the above argument, it represents a con
straint which does not change the form of the conditions of equilib
rium. This fact, immediately, leads to some physically interesting
conclusions: (1) The envelope can be laid in such a way as to cut off
a part of any of the phases. Therefore, the conditions of equilibrium
cannot depend on the total masses and volumes of the phases but only
on their specific properties like temperature, pressure, composition,
etc. (see below). (2) It can be laid so as to include only two of the
phases and exclude all the rest. Therefore, the equilibrium of two
phases is determined only by their own properties and is not affected
by the presence of other phases. (3) Since this is true for any pair of
phases, the problem of equilibrium of several of them can be reduced
to the simpler one of equilibrium between two. It also follows from
this that two phases which are in equilibrium with a third are in
equilibrium between them. (4) In general, a thermodynamical sys
tem is not completely defined unless it is known how it is delimited at
its boundary. However, it is clear from the preceding discussion that
any boundary conditions can be replaced, without loss of generality,
by rigid, adiabatic walls enclosing the system. In this case, the con
straints introduced by the walls are not only auxiliary but necessary
as forming part of the definition of the system. In the following sec
tions we shall use the expedient of the rigid, adiabatic envelope for all
these purposes.
A further example of an auxiliary constraint in geometrical guise
is a semipermeable partition or semipermeable membrane, interposed
between two phases, which lets through one kind of molecules but is
impenetrable to all other kinds. This device can be also regarded as
a part of the definition of the system, inasmuch as it is the thermo
dynamical criterion for the difference or identity of molecules: two
particles are different if it is possible to find a semipermeable membrane
102 TEXTBOOK OF THERMODYNAMICS VI 39
which lets through the one but holds back the other. 1 It is, there
fore, perfectly legitimate to introduce semi permeable partitions which
are selective not only with respect to chemical differences of mole
cules but also with respect to differences in the physical (quantum)
state of otherwise identical particles. 2
The notion of specific (also called intensive) properties used above
requires some amplification. It is easy to define it accurately for
phases which are described by the variables />, T, N\, N 2 , . . . N ff .
Any property which does not change when all the N k are increased in
the same proportion e (p and T remaining constant) is a specific prop
erty. For instance, the composition is best described in a specific way
by the mol fractions
oc k  N k /N, 
N = Nl + N 2 + . . . + N,. J (6 ' 8)
It follows from the definition of the mol fraction that
*i + *2 + . . . + x a = 1. (6.08')
In general, a function F n (N\, . . . NJ which changes in the pro
portion e n , when all the variables N k are increased in the proportion e,
F n (eNi, . . . eN.) = e"F n (Ni, . . . NJ, (6.09)
is called, in mathematics, a homogeneous function of the nth degree. It
satisfies Euler's equation
which is easily derived (compare exercise 62) from the definition (6.09).
Mathematically speaking, the specific quantities are, therefore, homo
geneous functions of the degree zero in the mol numbers Nk, since they
satisfy the condition
. . . NJ. (6.11)
Of course, F depends also on p and T, but we shall compare only
states and systems at the same temperature and pressure and need
*An interesting discussion of these matters will be found in a paper by
L. Szillard, Zs. Physik 32, p. 840, 1925.
1 Einstein pointed out the paradox that a semipermeable membrane must dif
ferentiate between two particles, be the difference in their states ever so little.
This difficulty disappears in the quantum theory where the states are discrete
(compare: J. Von Neumann, Math. Grundlagen d. Quantentheorie, p. 197.
Berlin 1932).
VI 39 GENERAL CONDITIONS OF EQUILIBRIUM 103
not bring these variables in evidence. The factor e may have any
value, and, in particular, we may put e = l/N, whence
Fo(Ni, ...#.) FQ(XI, . . . *.). (6.12)
In a specific quantity we may replace the mol numbers Nk by the mol
fractions Xk without changing it.
On the other hand, the volume F, the internal energy Z7, and the
entropy 5 are additive quantities. When all the mol numbers Nk are
increased in the proportion e, the mass of the whole system increases
in the same proportion, without change in its composition. The quan
tities F, U, S are proportional to the mass and also increase in the
proportion e (p and T remaining constant). The thermodynamic
potential is defined as $ = U TS + pV and has, obviously, the
same property
. . . N 9 ). (6.13)
The thermodynamic potential, like all the other additive functions
of thermodynamics, is a homogeneous function of the first degree. Hence
its derivative with respect to Nk, or the partial molal thermodynamic
potential, <pk = d$/dN k (compare section 35), is a homogeneous function
of the degree zero, as is easily shown by differentiating (6.13) with
respect t& Nk. Therefore, $ and ^* satisfy Euler's equation (6.10)
with n *= 1 and n = 0.
+ ...+
(614)
As 3^*/3^y = d$j/dNk, this can be also written
Every term of this expression is a homogeneous function of the
degree zero. We can, therefore, make use of the property (6.12) and
replace the mol numbers by the mol fractions
Although this relation was already given by Gibbs, it is usually
called Duhem's equation.
104 TEXTBOOK OF THERMODYNAMICS VI 40
Exercise 62. Prove Euler's eq. (6.10) from (6.09). (Directions. Differentiate
(6.09) with respect to e, denoting, for short, eN k  *. Then put e = 1; & = Nk)>
Exercise 63. Check by direct calculation the statement that Euler's equation
is correct for the homogeneous (inx t y,z) function F n  x a y*z n ~ a ~ b . Therefore, it
must, be true for any sum of such terms of the degree n.
40. The temperatures and pressures of phases in equilibrium.
Let us consider a heterogeneous system consisting of simple phases.
Its total internal energy U is the sum of the internal energies C7 (i) of the
individual phases
u = u m (6 * 17)
11
In the same way, the total entropy S is additive with respect to the
individual entropies S (i)
>, (6.18)
and we define as the total volume V the sum of the volumes F (0 of the
phases
a
**. (6.19)
We can choose 5 (i) and V (i) as the two thermodynamical variables
characterizing the state of the phase (i). It will, further, depend on a
number of parameters describing its chemical composition. In gen
eral, the phases will not be chemically pure substances but mixtures
of several or many chemically pure constituents or components. Sup
pose that they contain Ni mols of the first component, N<z of the
second, and so on, the number of components being <r. The internal
energy of the phase (i) is, then, a function of the following variables
. . #<*>), (6.20)
although a part of the mol numbers N may be equal to zero in some
of the phases. This is the form we have considered in section 31 : the
variation of the total internal energy U is, therefore, in accordance
with eqs. (5.06) and (5.07),
Applying auxiliary constraints, as explained in the preceding sec
tion, we shall at first suppose that the mol numbers do not change
VI 40 GENERAL CONDITIONS OF EQUILIBRIUM 105
(dN k (i} = 0) and shall find only the conditions controlling the thermo
dynamical variables.
The conditions of equilibrium are, according to eqs. (5.03) : dU = 0,
8S = 0. It is mathematically convenient to regard dU = as the
primary condition. It now takes the form
dU = > (T (i} 5S (i) ( '>$F (i) ) = 0. (6.22)
The condition
Of
55 = ^ SS (i) = (6.23)
<i
we regard as subsidiary. Another subsidiary restriction is obtained
if we replace the border conditions of the system by the assumption
that it is surrounded by a rigid adiabatic envelope (compare the pre
ceding section) : We have then F = const, or
In the case when the several phases are separated by natural sur
faces of discontinuity and each of them can expand, virtually, at the
expense of the others, the variations <5S (i) , 5V (i) are not subject to any
other restrictions so that eqs. (6.22), (6.23), and (6.24) exhaust the
problem. Multiplying eqs. (6.23) and (6.24), respectively, by \i and
\2 and adding them to (6.22)
a
5
= 0, (6.25)
11
where the dS (i) and 6F (i> must be considered as independent (compare
section 38):
The temperatures of all the phases are equal to the same function
which has, therefore, the physical meaning of the uniform temperature
T of the whole system (Xi = T). In the same way, the pressures
of all the phases are equal to p
p M =p. (6.26)
In short, the temperature and pressure in a system in equilibrium are
uniform. However, this applies only to the case when each of the
phases can change its volume. In the opposite case, when the con
106 TEXTBOOK OF THERMODYNAMICS VI 40
straints are such that the volume of each phase remains unchanged
(y> = const.), the condition (6.24) must be replaced by a separate
conditions
57>=0. (6.27)
This will occur, for instance, when each of the phases is enclosed in
rigid heatconducting walls, but also in some cases which are less
trivial, as we shall see in sections 107 and 127. In applying the
method of Langrangean multipliers, we take eq. (6.23) with the factor
Ao and (6.27) with the factors X (i) and add them to eq. (6.12)
S[(r>
= o, (6.28)
whence
r =  Ao = T, />> = A (i) . (6.29)
The temperature is uniform throughout, as in the previous case, but
the pressures of the phases are all different. It is quite obvious how
these considerations are to be extended to the mixed case when the
volume of some of the phases is constant while others can expand.
The case when some of the phases are adiabatically enclosed and their
entropies are constant (dS (i) = 0) is also possible but hardly worth our
consideration, since it is one of dynamical, rather than thermodynami
cal, equilibrium.
We have obtained a partial solution of the problem of equilibrium,
with respect to the thermodynamical variables T and />, and can now
turn to the other side of the problem relating to the virtual changes
of the mol numbers. In order to deal with it, we shall again impose
upon the system suitable constraints. We already know that, in
equilibrium, the temperature T is uniform while the pressures p (i) of
the phases may or may not be equal. We shall now consider only
virtual changes in which dT = and dp (i) = 0. It was shown in sec
tion 36 that, under these circumstances (when the temperature and
the local pressures are kept constant), the function <f> tends to decrease
and equilibrium is possible only when it has reached its minimum
(&typ. T = 0. (6.30)
The total thermodynamic potential is represented by the sum over
the phases
VI 41 GENERAL CONDITIONS OF EQUILIBRIUM 107
In forming the variation T and p are to be kept constant, whence
In general, the virtual changes 5N t (i) are not independent but sub*
ject to subsidiary conditions determined by the nature of the system.
Exercise 64. Consider eqs. (6.22), (6.23), and (6.24) for the case of only two
phases and solve them without use of Lagranean multipliers. Show that this
method can be used to obtain the general result by the reasoning of section 39.
Exercise 65. Take the case of p and T being uniform throughout the system
and generalize the argument of section 32 which led to the formula (5.08) for systems
with several phases. Show that this is an alternative way of obtaining the condition
(6.31).
41. Number of phases in equilibrium. The phase rule. In order
to derive the explicit equations of equilibrium from the condition
(6.31), it is necessary to eliminate from it the dependent virtual
changes dN k (l) by means of the subsidiary conditions. It was pointed
out in section 38 that this can be done either before taking the varia
tion (6.30) of the function <J>, or after. The former procedure consists
in expressing a part of the arguments N k (i) of the function $ by the
remainder and so reducing their number as far as possible. We say,
then, that the thermodynamical potential is expressed in terms of the
mol numbers of the independent components of the system. To illus
trate this by an example, let us take the case that N\ mols of hydrogen
(H2) and N% mols of iodine (12) are brought into the same vessel.
The two gases will partially react and form a certain amount (N% mols)
of hydrogen iodide which will be the third constituent of the mixture
but not an independent one: in the state of equilibrium N& is com
pletely determined by N\ and N%, so that hydrogen and iodine can be
taken here as the independent components. We see, moreover, from
this example that the method is of no help in the treatment of con
crete cases, because we must have solved the problem of equilibrium
and found the expression of Na, in terms of NI and N2, before we can
eliminate it from the thermodynamical potential. This is, in fact,
characteristic of all thermodynamical systems: the main part of the
subsidiary conditions between the virtual changes 6N k (i) are given in
the variational form (6.06), and not in the integral (6.03), and often
there is no practicable way of eliminating the dependent mol numbers
N k (i) except by means of the final solution of the problem. However,
some interesting conclusions can be drawn from the theoretical possi
bility of this elimination: we simply suppose that the system is reduced
108 TEXTBOOK OF THERMODYNAMICS VI 41
to its independent components, without attempting to carry out this
reduction in any special case.
Generally, those constituents of a system should be chosen as its inde
pendent components which cannot be converted into one another by any
reaction going on in it. In the above example hydrogen and iodine
satisfy this requirement, but, in general, the number of independent
components of a system is, by no means, identical with the number
of chemical elements of which it consists. For instance, if all the
phases of a system consist of water in different states of aggregation
(ice, steam, liquid water), hydrogen and oxygen are not independent
components, because the mol number of the one is defined by that
of the other, viz. stands to it in the constant ratio 2:1. We have
here a system of only one independent component. Strictly speaking,
the number of components depends on the accuracy with which we
wish to describe the system : there is always a certain amount of disso
ciation in steam, and, because of the different solubility of hydrogen
and oxygen, the phases will not contain exactly equal mol numbers
of the two elements. However, these effects are entirely negligible
at ordinary temperatures. An example of a system in which the num
ber of independent components is larger than the number of elements
is presented by a mixture of hydrocarbons. There are many hydro
carbons which do not react under ordinary conditions, so that the mol
numbers of hydrogen and carbon are quite insufficient for a description
of such a system. The ultimate criterion for the choice is, of course,
that a virtual change of the mol number in one of the independent com
ponents should not necessitate any virtual changes in the others.
The notion of independent components is sufficiently clear from
this discussion. Suppose that their number is ft and that the mol
numbers of all the constituents, the independent as well as the depen
dent, are eliminated from the functions <> (i) and replaced by the masses
Mi (i \ M2 (i \ . . . Af e (i) of the independent components in the several
phases. Equation (6.31) becomes, then (if we drop the subscripts
(6 ' 32)
The only subsidiary conditions which remain are now those which
apply to each independent component separately. We shall restrict
ourselves here to the case that the phases are in direct contact, so that
a component can pass freely from one phase into another. As was
pointed out in the preceding section, the border conditions of the
system as a whole can be replaced by an imaginary rigid envelope
VI 41 GENERAL CONDITIONS OF EQUILIBRIUM 109
surrounding it. This implies that the total mass of each component
a _
must remain constant. /, Af* (0 = const, or taking the variation,
0, (* = 1, 2, . . . ft).
Multiplying these equations, respectively, by the Lagrangean mul
tipliers X* and adding them to (6.32),
\
). ft
 +X*
where the virtual changes BM k are completely independent. There
follows the system of equations
=  X*. (6.33)
The index k goes here from 1 to ft, so that this formula represents
aft relations. If we eliminate from them the functions X*, there
remain Zeq = aft ft equations between the partials of the functions
3> (i) . What is, on the other hand, the number of variables on which
these partials depend? It was shown in section 39 that the partials
of <i> (i) , with respect to N k (i \ represent specific quantities. The reason
ing used in that demonstration is in no way affected, if the masses
M k (i) are regarded as the arguments of these functions, instead of the
mol numbers. The partials 9* (i) /3^ (0 are, therefore, homogeneous
functions of the degree zero in the masses. In analogy to the mol
fractions we introduce the mass fractions
(6.34)
and conclude from eq. (6.12) that 3* ( * ) /9Af* (0 can be represented in
terms of the variables y k (i) . Only ft 1 of these fractions are inde
pendent because there exists the identical relation
The number of variables defining the composition of the phase (i) is,
just ft 1, and the number of them in all the a phases is aft a. To
this must be added the two thermodynamical variables, the common
temperature T and pressure p of all the phases. The total number of
110 TEXTBOOK OF THERMODYNAMICS VI 41
variables by which the partial derivatives of the system (6.33) are
determined is, therefore, Z var = a/3 a + 2.
A system of equations is only then compatible and capable of
being satisfied when their number does not exceed the number of
variables on which they depend (Z eq g Z var ). This condition restricts
the number of phases which can be simultaneously in equilibrium:
from a/3 (I afl a + 2 there follows
a ft + 2. (6.36)
This inequality expresses the famous phase rule of Gibbs: The
number of phases coexisting in equilibrium cannot exceed the number of
independent components by more than two.
Let us consider the main cases that can arise. Examples of them
will be treated in the following chapters. The maximum which a can
reach is
a = + 2, (6.37)
it will be attained when the number of variables is just equal to the
number of equations. All the variables are then completely deter
mined, the equilibrium can exist only at one definite temperature and
one definite pressure, and the composition of all the phases is also
quite definite. This state is often called the fundamental state of the
system.
In the case
a = /3 + 1 (6.38)
the number of the variables exceeds that of the equations just by one:
the value of one of the variables can be chosen arbitrarily, and this
fixes the values of all the others. For instance, if we specify the tem
perature, the pressure and the compositions are completely deter
mined by it, etc.
Similarly, when
a = 0, (6.39)
two of the variables can be assigned arbitrary values: if p and T are
specified, the compositions of the phases are definite.
Finally, in the cases a < /3, the compositions of the phases are not
yet defined by the temperature and pressure. There are, then, so
many cases of equilibrium possible that hardly any rules of general
validity can be given.
In conclusion, it will be well to emphasize that the phase rule deals
only with the number of phases in equilibrium. We have mentioned
VI 42 GENERAL CONDITIONS OF EQUILIBRIUM 111
water in its different states of aggregation as an example of a system
with one independent component (ft = 1). According to the phase
rule not more than 1+2=3 phases of it can coexist (viz. vapor,
liquid water, ice), and the simultaneous equilibrium of all three is
possible only in the fundamental state at a definite temperature and
pressure (compare also section 48). However, everybody knows that
he can throw a lump of ice into a bucket of water and that there will
be, at all temperatures, water vapor on top of the system. In this
case, the three phases are not in equilibrium, and, in the long run, one
of them will disappear, but the coexistence without equilibrium may
last hours. In applying the phase rule one must be quite sure that
equilibrium is, actually, established.
42. Explicit equilibrium conditions. We return now to the
description of the phases by the mol numbers of all their chemically
distinct constituents and not only of the independent components.
We shall make use of the notation (5.38) defining the partial molal
thermodynamic potential
>
(6  40)
It will be well to recall here that it is a generalization of the ordinary
molal potential. In the special case, when the phase is chemically
pure, there is only one constituent, 3> (i) = N (i) <p (i \ where <p (i) is inde
pendent of N (i \ Then ^ (i) becomes identical with <p (i) .
The condition (6.31) now takes the form
' = 0, (6.41)
and is valid for all possible virtual changes dN k (i) consistent with the
nature of the system.
The virtual processes in the system can be reduced to two simple
types. In the first place, a small amount of the constituent k can pass
from a phase (1) into an adjacent phase (2). The molecular weights
MA: (I) and jufc <2) of this component in the two phases need not necessarily
be the same because there is the possibility of association. The mass
added to the first phase is, then, M* (1) &W 1} and that added to the
second n k (2) 5N k (2} . These masses must be oppositely equal
or
W> : 6N k   1 : ^  *> : ,*>. (6.42)
(1) (2)
112 TEXTBOOK OF THERMODYNAMICS VI 42
The symbols v k are integral numbers with the following meaning.
Suppose that j>* (1) mols of the component k are removed from phase
(1): they are converted into v k (2} mols belonging to phase (2). If no
association takes place in either phase, we can put 1>* (1) = v k (2) = 1.
There are no conditions in the system which should prevent us
from putting all the other virtual changes bN k equal to naught, in the
way of an auxiliary constraint. Therefore the condition (6.41) reduces
to
A$ SE v k Tp k + v h *W> = 0. (6.43)
A<i> can be interpreted as the change of the total thermodynamic
potential which takes place in the finite virtual process of reversibly
transferring v k m mols of the component k from the first phase into
the second (where they are converted into v k (2) mols). The temperature
and pressure are supposed to remain constant, maintained so by a
suitable arrangement of heat baths and loads. Of course, the process
will be attended by the system receiving (positive or negative) heat
from the heat baths and doing work against the loads.
Sometimes it is convenient to express the composition not in terms
of the mol numbers N k (i) but in terms of the masses M k (i) = n k (i} N k (i) of
the individual components. We shall denote the partial specific
thermodynamic potentials (referred to unit mass)
Because of (6.42), /** (1 W n = Mt (2 V 2) , and (6.43) reduces to
?Af* (1>  ?W 2) = 0. (6.45)
The second simple type of process we have to consider is a chemical
reaction between several constituents of the same phase. Let us take,
as an example, the equation of the reaction between hydrogen and
iodine
2HI = H 2 + I 2 ,
or
2HI  H 2  I 2 = 0.
Like the equations of all chemical reactions it is of the general type
viGi + v 2 G 2 + . + v 9 G 9 = 0, (6.46)
where GI, . . . G are the chemical symbols of the different substances
(components) an4 ?i, . . . v 9 the numbers of mols with which they
take part in the reaction. We may observe here that the process of
VI 42 GENERAL CONDITIONS OF EQUILIBRIUM 113
transferring matter from one phase into another, as considered above,
can be represented by the same sort of a symbolical equation
>=0. (6.47)
Both equations are included in the general form
> = 0. (6.48)
We shall usually consider a definite direction of the reaction, namely,
that in which it is endothermic and the heat of reaction positive. Con
sequently, we shall attribute positive signs to the numbers vk of those
components which are produced, and negative to those which are
consumed, in the endothermic reaction.
When eq. (6.41) refers to a process of the type (6.46), the corre
sponding virtual changes dNk must be taken proportional to the num
vers vk
v%: . . . : v ff .
(The numbers vk and variations dNk for the components not taking
part in the reaction can be put equal to zero because of the principle
of auxiliary constraints). The condition becomes
A$ E= vjQt + . . . + Vj p 9 = 0. (6.49)
As in the preceding case A<i> must be interpreted as the total change
of the thermodynamic potential in the finite virtual process corre
sponding to the transformation of *>i, v v mols of the several com
ponents.
Equations (6.43) and (6.49) represent partial solutions of the prob
lem of equilibrium. The complete solution is given by the simulta
neous system of equations corresponding to all processes of these two
types possible in the thermodynamical system. Both these equations
are included in the general form
in which a part of the coefficients v k (i) are equal to those of (6.43) or
(6.49) and the rest vanish. It will be useful to find the partial deriva
tives of A< with respect to the pressure and temperature: from
eqs. (5.37)
114 TEXTBOOK OF THERMODYNAMICS VI 42
and (3A$/9r) p = AS, where AFand AS represent the changes of the
total volume and the total entropy in the process. This process being
isothermal (as well as isobaric), we can apply to AS the formula (4.11),
AS = Qp/T, where Q p is the isobaric heat of the reaction
(6.52)
Moreover, taking into account the condition A3> = 0, we can write
Qp (6.53)
Exercise 66. What is the structure of A 7 and Q p in terms of partial molal
quantities? Show that
, Q,
*
CHAPTER VII
PHASES OF A CHEMICALLY PURE SUBSTANCE
43. Equilibrium of two phases. If all the phases of a system con
sist of the same chemically pure substance, in different physical states,
it contains, according to section 41, only one independent component
(ft = 1). The phase rule (6.36) tells us, then, that no more than three
phases can be simultaneously in equilibrium. There exist substances
which are known in more than three states. For instance, sulfur
occurs in the vaporized, liquid, monoclinic crystalline, and rhombic
crystalline states, but only three of them can coexist in equilibrium.
The problem of thermodynamics in relation to pure substances is,
therefore, reduced to the consideration of the equilibrium of one phase,
two phases, or three phases. The properties of a single phase were
sufficiently discussed in Chapters I to IV, and we shall restrict ourselves
here to the cases of two and three phases, starting with the twophase
system as the simpler problem.
The only possible virtual process in such a system is that repre
sented symbolically by eq. (6.43): v (1) G (1) + v (2) G (2) =0, viz. the
transformation of v (2} mols of the second phase into v (l) of the first.
It will be more convenient to define the numbers v (1) and v* both as
positive and to indicate the fact that the second phase is consumed in
the process by writing its equation as v (l) G (l) j/ (2) G (2) with
instead of (6.42). It was pointed out in the preceding section that, for
a pure substance, the partial thermodynamic potential and the ordi
nary molal are identical ^ = <p. The eq. (6.31) of equilibrium takes,
therefore, the form
A* = (1 V (1)  v (2 V 2 >, (7.02)
or, in the special case when there is no association and the molecule
is the same in both phases (> (1) = v (2) = 1),
<<>  > (2) . (7.03)
In words, the molal thermodynamical potentials in two adjacent
115
116 TEXTBOOK OF THERMODYNAMICS VII 43
phases are equal. The molecular constitution of the condensed phases
is not always known and, in some cases, even not well defined. For
this reason, many authors refer the quantities <p (i) of all the phases to
1 mol of the vapor. In particular, this is unavoidable in the chemical
literature which has to present results of measurement on all sorts of
substances, including those whose molecular structure is not yet suf
ficiently investigated. The purpose of this book is, however, a different
one: we are not so much interested in the presentation of a large
amount of experimental material as in its theoretical penetration.
The questions of association and dissociation play an important role
in the interpretation of the experimental behavior of substances, and
they can be conveniently studied if the fundamental equation of
equilibrium is written in the general form (7.02). We shall, therefore,
make this formula the basis of the following discussions (compare also
section 60).
In general, the coefficients *> (i) are integers, completely determined
by the nature of the system, while the functions <p (i) depend only on
the common temperature and pressure of the two phases, p and T
are, therefore, the only variables which enter into eq. (7.02) or (7.03),
and, if one of them is arbitrarily chosen, the other is determined by
this relation. It was pointed out in section 41 that this is always the
case when the number of phases is larger by one than the number of
independent components (a = ft + 1). When the equilibrium of the
liquid and the gaseous phase is considered, the temperature which
corresponds to a given pressure, in virtue of the condition (7.02), is
called the boiling temperature or the boiling point. Vice versa, if the
temperature is regarded as given, the pressure corresponding to it is
known as the boiling pressure. Similarly, one speaks, in the case of the
coexistence of the liquid and the solid phase, of the freezing point and
the freezing pressure. More general terms including all cases of a pair
of pure phases are transformation temperature and transformation
pressure.
Let AC and DF in Fig. 12 represent, respectively, the isothermals
of a pure substance in its liquid and gaseous states. Suppose we have
a single phase of it corresponding to the point A of the diagram.
Keeping the temperature constant we increase the volume reversibly,
thus allowing the pressure slowly to drop: the representative point
of the system will move along the isothermal towards C. Let the
point B correspond to the boiling pressure for this isothermal at which
the coexistence of the liquid and the vapor phase is possible. The
question arises now: will the substance, after reaching this point,
remain in a single liquid phase and continue to move along the branch
VII 44 PHASES OF A CHEMICALLY PURE SUBSTANCE
117
B, or will it begin to vaporize at constant pressure and be converted
into a twophase system whose states are represented by the horizontal
BE? What is the relative stability of the states M and N lying
on a vertical and corresponding to the same volume? The results of
the preceding chapter are insufficient to solve this problem because
they regard the number of phases as given: the criteria given there
are satisfied in both cases. To obtain an answer, we must fall back
on the fundamental principles governing thermodynamical equi
librium as expounded in Chapter V. Since the processes BM and BN
are both isothermal and lead to the same
end volume, we have to use the criterion
adapted to this case (section 36): The
work function ^ tends to decrease, and,
of two states of the same temper
ature and volume, that with the lower
is more stable. Let us now evaluate
the work functions &M and ^N in the
states M and N by comparing them
with ty B in the state B. We make use
of eq. (5.22) which applies to any iso
thermal process: ^B ^M = WBM a
the right sides of the equations repr
system, respectively, in the processes
find Vtr VM = WBM WBN. In
represented by the areas under the
. Isothermals in equi
lum of liquid and vapor.
M lies lower than N, we have
^fN < ^M The states on the h<
work function and of higher stabi
that the states represented by
impossible but sometimes occ
(supercooling). However, they
and the system is always read' {
This reasoning can easily be
transformation pressure or
divides into two phases. J
tional conditions.
:ur
cor;
it is
VB ^N = WBN, where
. the work done by the
and BN. Subtracting we
iiagram WBM and WBN are
and BN\ and, since
We find, therefore,
E are the states of lower
<?have mentioned in section 5
anch BC are not absolutely
' exceptional circumstances
d to an unstable equilibrium,
\ver into the twophase state,
o all analogous cases: at the
i a pure substance, generally,
5 a single phase only in excep
Exercise 67. When the
and vapor), there are two ;
represented by the curve an^
prove that the two shaded , ___
* areas
44. The
auy
be directly used for the call u ' at * on
V
^ o /fc . equation applies (continuity of liquid
,O/ cesses between B and E (Fig. 1, p. 10)
n ' , line. Applying eq. (5.21) to both of them,
1 are equal.
equation. Equation (7.02) cannot
transformation temperatures
118
TEXTBOOK OF THERMODYNAMICS
VII 44
because the explicit form of the functions <p (i) is known in only very
few cases. However, its differential form gives a relation between
easily measurable quantities. Let us consider the equilibrium of our
twophase system in two cases: at the temperatures and pressures
r, p and T + dT, p + dp. The respective equations (7.02) are
A$>(r, p) = and A<i>(r + dT, p + dp) = 0. According to the
definition of the differential of a function, the difference between these
two equations is
v (*)+&).* <"*>
We can sX^';iite the expressions (6.51) and (6.52) for the partial
derivatives. heat of the process (Q p ) is, in this case, the latent
heat of trans *ion L (compare section 10). We find, therefore,
. V,
\
This relation was fc
the discovery of the first
from the modern point o*
called the ClapeyronClau.
transformed in the process
we shall denote the latent he
AFj (v (1) i; (1) j/< 2) z; (2) )/j/<
and before the transformat
(7.05)
^ '
"* ^French engineer Clapeyron before
aTermodynamics (1832). Its deduction
^ ^ due to Clausius. It is, therefore,
& ^ion. If the mass of the material
^ e ^s to 1 mol of the vapor phase,
^ , hile the change of volume becomes
"Terence of the molal volumes after
' referred to 1 mol of the vapor)
(7  06)
This formula has been ext
to be accurate within the limi* ve ^ '
we may take water at the no sf ex ?? n
of volume is AF 2  30 186 c/ boiling
mm/deg. The conversion lo1 .
ested by experiment and found
'mental error. As an example
point (373. 1): the change
T lg Q1S ^ dp/dT = 2? 12
e . mm/deg = (1.01325/760
X 4.1852 X 10) cal deg" 1 cm
' 3 *
The formula (7.06) gives, there
direct measurement gives / = 972v, . t
We give in Table 8 the normal ^ *** C
melting points (7 of the more
together with the latent heats of
at these temperatures.
X 10 5 cal deg 1 cm" 3 .
'6 cal mol 1 , while the
L Ailing points (T B ) and
e  c mentary substances
and of fusion (W
VII 44 PHASES OF A CHEMICALLY PURE SUBSTANCE
TABLE 8
119
Element
T B
IB (cal mor 1 )
IB/TB
TF
IF (cal mol 1 )
He
4.2
24.5
5.7
Ne
27.3
416
15.3
24.5
57
A
87.5
1 540
17.6
84.0
267
Kr
121.4
2240
18.4
104
360
Xe
164.1
3200
19.5
133
490
H 2
20.5
220
10.7
14.0
28.4
N 2
77.3
1370
17.7
63.3
170
2
90.2
1640
18.2
54.8
105
8
161
3570
22.2
22
C1 2
238.6
4900
?0.5
171.6
1620
Br 2
332.0
7150 20.9
266.0
2560
I 2
457.5
105'
23.0
386.7
4000
Na
1153
25
21.8
370.7
633
K
1033
20
19.8
335.5
518
Cu
2570
11
43.4
1356
2750
Ag
2220
)
26.8
1233.7
2630
Au
2870
30.6
1336
3180
Zn
1180
20.7
692.6
1690
Cd
1040
JO
24.5
594.1
. 1480
Hg
630.1
)0
23.0
234.3
556
Al
2070
26.0
933.2
1910
Si
2870
)
14.3
1690
Sn
2530
I, <
30.6
505.0
1670
Pb
1890
4610v,
24.4
600.7
1120
Bi
1720
46100
26.8
544
2600
S
717.8
2190
30.5
392.2
262
W
6170
223000 36.2
3643
Fe
3270
90700 28.6
1808
2670
Pt
4570
167000 36.6
2028
5250
.J
In the case of the transformations solid liquid and liquid
* vapor, the latent heat L is positive. The volume of the vapor is,
moreover, always larger than the volume of the same mass of liquid,
AF> 0, so that the boiling temperature rises with the pressure
(dTB/dp > 0). For instance, in the case of water we have the follow
ing conditions:
TABLE 9
t
PS
6.025 X 10
50 /'
O.fi'17
100
1.0000
120
1.9594
200
15.341
370.4 C
217.72atm
On the other hand, the specific volume of the liquid phase may be
either larger or smaller than that of the solid in equilibrium with it.
120
TEXTBOOK OF THERMODYNAMICS
VII 45
For instance water (at C and 1 atm) has a 9% smaller volume than
ordinary ice. Therefore, AF and dT/dp are in this case negative, as
was pointed out by James Thomson (elder brother of Lord Kelvin).
The melting point falls, therefore, with increasing pressure: when
ice at a temperature slightly below C is subjected to pressure, it is
brought into a state in which its T is above the melting point. It is,
therefore, converted into water but freezes again as soon as the
pressure is released. This phenomenon is called regelation, and it
accounts for the plasticity of ice which permits, among other things,
the flow of glaciers. The ease of skating on ice is also usually attrib
uted to regelation : enough water ir' '.elted under the pressure of the
sharp edges of the skates to k*> /
In the case of vaporization, c
the formula (7.06) can be simplih
so low that the equation of state (
sufficient approximation, to the gasv
these conditions the molal volume .
neglected compared with V2. In fact,
ratio V2/vi for water (e.g.) is 1676, and
law of perfect gases can be used, it is
obtain an approximate form of eq. (7.C
/lubricated.
are sometimes such that
0^ the vapor pressure is
perfect gases applies, with
ihase: V2 = RT/p. Under
the liquid phase can be
at 100 C and 1 atm, the
* temperatures, when the
arger. In this way, we
_
RT*
which is useful for many purpose*
Exercise 68. Calculate the latent / of water, at 100 C and 760 mm Hg,
from eq. (7.07). Use the value dp/d /> 27.12 mm/deg. Compare the results
with the exact value as given in text.
Exercise 69. Calculate the latent heat of mercury at its normal boiling point
(tB  357.3, p  760 mm). Use the value dp/dt = 13.81 mm/deg and eq. (7.07).
Exercise 70. Calculate the coefficient dT/dp for the fusion of ice at C and
1 atm, using the following data (per 1 g of the substance): l g = 80 cal g" 1 ,
v\ = 1.000 cm 8 g" 1 , j = 1.091 cm 8 g" 1 . Apply suitable conversion factor to p.
Exercise 71. Naphthalene melts at 80. 1 C, its latent heat of fusion is 4563
calories per mol, and the increase in volume on fusion is 18.7 cm 8 mol" 1 . Find the
change of the melting point with pressure.
Exercise 72. The transition point of mercuric iodide, red > yellow, is at
127 .4 C. The latent heat absorbed in the change is .150 cal mol" 1 , and the approxi
mate change of volume 5.4 cm 8 mol" 1 . Find dT/dp for the transition point.
45. Temperature dependence of the latent heat Another rela
tion between measurable quantities can be obtained by differentiating
VII 46 PHASES OF A CHEMICALLY PURE SUBSTANCE 121
eq. (3.23) of section 17, dL = d(X (1)  X (2) ) = </(*> (1) x (1)  ^ (2) x (2> ) and
substituting the expressions (4.47) for dx We find
dL = AC p <*r
where AC P and AF denote the changes of capacity and volume in the
transformation :
AC P = v (1 V 1}  v (2 V 2) > AF  iW  / 2 V 2) . (7.08)
Dividing the equation by dT and replacing dp/dT by (7.06)
In the special case of vaporization at comparatively low pressure
we can again make the simplification of neglecting v (2) (of the con
densed phase) and applying to v (1) the equation of perfect gases:
log AF/3r) p = 1/r. The relation is then reduced to
 *Cr (7.10)
Exercise 73. Calculate dl/dT (per 1 mol = 18.0) for boiling water (at 100 C,
1 atm) with the help of the following data: in liquid water c pz 18.0 cal, in vapor
c Pl = 8.8 cal (both per 18 g, at 100 C), AF f = 30 186 cm 3 mol 1 , (QAKj/aDp =
84.66, I = 9720 cal mol" 1 . Compare the results given by the two formulas (7.09)
and (7.10). (The measured value is 11.0).
46. Approximate and empirical expressions for the transforma
tion pressure. The temperature dependence of the latent heat is,
usually, slight. Especially, in the case of vaporization AC P in the for
mula (7.10) is small compared with L. As a first rough approximation
one can, therefore, put L = LQ = const, or / = /o = const. Sub
stituted into eq. (7.06) this gives for the pressure of vaporization
(7.11)
(B being a constant), a formula due to Van der Waals, who wrote it
(for liquids) in the form
(7.12)
which for a number of substances gives an approximate representation
of the measurements even at high pressures. We shall return to this
equation in the next section.
122
TEXTBOOK OF THERMODYNAMICS
VII 46
It has been found since that the formula (7.11) is insufficient for
an accurate representation of the vapor pressure of liquids but can be
used for that of crystalline solids. Even in this case better approxima
tions are obtained by substituting into (7.07) the expansion
/ = / + 1 V T + faT 2 + . . . . For instance the vapor pressure of ice
between 90 and C is represented by l
944.C c
+ 8.2312
 0.0167006r
+ 1.205 X 10~ 5 r 2  6.757169. (7.13)
As to the vapor pressure of liquids, the very accurate modern
measurements cannot be adequately represented even by formulas
with many terms.
For the dependence of the pressure of fusion upon temperature a
theoretical formula has not yet been advanced, and one has to rely
on empirical expressions. According to Simon and Glatzel 2 the follow
ing formula gives satisfactory results:
logio(/> + a)=c logio T F + 6. (7.14)
They give the following values for the constants (Table 10), pro
vided the pressure is measured in kg/cm 2 = 0.9678 atm.
TABLE 10
Substance
a
b
c
Helium*
17
1 . 5544
1.236
Sodium
12400
5.032
3.55
Potassium
3900
8.227
4.68
Rubidium
4 100
5.59
3.65
Cesium
2450
8.38
4.75
Mercury
45000
1.11
2.40
CO 2
4000
2.564
2.64
ecu
3 100
1.498
2.08
SiCU
5 100
0.459
1.39
Benzene
3 900
2.496
2.49
Chlorobenzene
5 100
1,970
2.41
Bromobenzene
5000
2.070
2.42
Nitrobenzene
6000
1.037
1.97
Chloroform . . ...
6000
0.466
1.83
Bromoform
5800
0.984
1.94
Aniline
5500
1,913
2.33
* The data for helium are taken from Simon, Ruhemann, and Edwards, Zs. phys. Chemie
(B) 2, p. 430, 1929.
1 E. W. Washburn, Monthly Weather Review 52, p. 488, 1924.
* F. Simon and A. Glatzel, Zs. anorgan. Chemie 178, p. 309, 1929.
VII 47 PHASES OF A CHEMICALLY PURE SUBSTANCE 123
Exercise 74. Calculate the latent heat / of sublimation of ice and the differ
ence Ac p of specific heats (between 1 mol of vapor and 1 mol of solid), at C. Use
eqs. (7.13), (7.07), and (7.10).
Exercise 75. At its normal point of fusion (38.9 C) mercury has the molal
volumes vi = 14.65 (liquid) and v, = 13.90 (solid). Its latent heat is / = 556 cai
mol" 1 . Check whether the formula of Simon and Glatzel is in satisfactory agree
ment with (7.05).
47. Vaporization as a corresponding state. It was stated in
section 27 that certain groups of substances approximately satisfy
what was called there the extended law of corresponding states. It is
meant by this that within such a group the quantities (u o)/r, 5
and pv/T are universal functions of the reduced variables. Since the
molal thermodynamic potential is defined as <f> = u Ts + pv, the
quantity (<f> uo)/T must also be expressible as
<p UQ . ^ /T i r\
^ =/(T,T), (7.15)
where / is universal for the whole group. The reduced volume does
not appear as a variable, because it can be eliminated by means of the
equation of state.
It must be assumed that the equations of state apply equally to
the gaseous and to the liquid phase, bringing out their continuity, in
the sense of section 5. Therefore, the formula (7.15) is also valid for
both phases. It follows then from (7.03) that, for substances which
do not associate, the boiling condition is determined by the universal
equation
/d)( 7r , T ) =/O>(T,T). (7.16)
Substances boiling at corresponding pressures have corresponding
boiling points. In practice the boiling point changes but little with
pressure. For instance, we see from the data for water on p. 119 that,
while the pressure is increased from 1 atm to 2 atm, the temperature
changes only by 20 or about 5%. For all the liquids satisfying
eq. (7.16) there obtain, in this respect, similar conditions. In fact,
it follows from it that the ratio
dT . dP _ dir .
~T ' p ~~ IT ' T
must be universal. This explains the fact noticed already by Guldberg 1
that many liquids have roughly the same reduced boiling temperatures
1 C. M. Guldberg, Zs. phys. Chem. 5, p. 374, 1890.
124
TEXTBOOK OF THERMODYNAMICS
VII 47
even when they all boil at the pressure of 1 atm, as appears from
Table 11.
TABLE 11
Substance
T c
T B
TB/T C
Ether
467
308
66
Ethyl acetate
523
346
66
Ethylene chloride
561
357
64
Alcohol
516
351
68
Benzene
562
353
63
Hydrochloric acid ....
Oxygen .
325
155
238
90
0.73
58
Carbon disulfide
Ammonia
545
404
319
234
0.59
58
Ethylene
282
163
58
Methane
191
109
57
Sulfuretted hydrogen.
Nitrous oxide
373
309
211
183
0.57
59
Stannic chloride
592
387
65
Water
637
373
59
Acetic acid
595
391
66
According to (4.11), the quantity l/T B represents the entropy
increase on vaporization
= s<i>  s< 2 >.
It is an entropy change between corresponding states and must be,
therefore, a universal function of TT, r according to (4.41). There
follows that I/TB must be the same for substances which boil at
corresponding pressures. However, both / and T B change but little
with the pressure so that, in practice, l/T B measured at any pressure
of the order of 1 atm, is roughly the same for many different sub
stances (Table 12). This fact is known as the rule of Deprez and
Trouton.
The table contains of course a group of selected substances. If
picked at random, the variations of Trouton's ratio are larger. Yet a
very marked deviation of the ratio IB/TB from the average value is,
usually, traceable to association. For instance, in acetic acid
IB/TB = 14.8, and this is due to the abnormally high molecular weight
of the vapor (97 instead of 60 corresponding to 2^02). Ethyl alcohol
shows the opposite deviation, IB/TB = 26.8, due to a normal vapor
but partly associated liquid. Equation (7.12) of Van der Waals is
VII 48
PHASES OF A CHEMICALLY PURE SUBSTANCE
TABLE 12
125
Substance
t B C
IB (cal)
h/T B
Benzene
394.8
13650
20.8
Chloroform
247
10900
21.0
Aniline
434
14920
21.1
Bromobenzene
155 9
9350
21.0
Acetal
102.9
7810
20.7
Toluene
109.6
7980
20.8
Sulfuretted hydrogen.
Stannic chloride
61.4
112
4490
7 900
21.2
20.5
Ethyl bromide
38.4
6500
20.9
Ethane
71 2
7 120
20.6
only a special approximate form of the general equilibrium condition
between a liquid and its vapor. To the extent to which the reduced
equation (7.16) is valid, the special form can claim to have universal
validity. In fact Van der Waals expected the coefficient a to
be a universal constant (for all substances). However, two things
must be borne in mind. In the first place, (7.12) is only a rough
approximation and a is, in reality, not a constant but a function of
temperature. In the second place, we should expect, at best, a uni
formity of a, not generally, but only within the groups of substances
satisfying the law of corresponding states. The analysis of the experi
mental material shows that the substances of such a group can be
arranged in a series so that a is nearly the same for any two successive
members of the series, but it changes, systematically and considerably,
from one end of the series to the other.
The emphasis on Trouton's rule, as on the law of correspondence
in general, has changed in the course of time. Soon after the discovery
of these regularities their importance and accuracy were being over
rated. Then the pendulum swung to the other extreme: the insistence
was altogether on the lack of rigor of these laws and on the numerous
exceptions from them. However, as long as it is borne in mind that
these rules are only rough approximations of limited range, they are
interesting and useful. For instance, deviations from Trouton's rule
are helpful in indicating association, as was pointed out above.
48. The fundamental or triple point. There remains to say a few
words about the third case mentioned in the beginning of this chapter.
It follows from the general considerations of section 41 that the
coexistence of three phases is possible only at a definite temperature
and pressure which jointly define the fundamental or triple point. In
126
TEXTBOOK OF THERMODYNAMICS
VII 48
fact, if we label the three phases by the numbers (1), (2), (3), we can
apply the results of the preceding sections to each pair of them: For
the equilibrium between (1) and (2), on one hand, and between (1)
and (3), on the other, we have the two equations of equilibrium of
the type (7.02)
the third y (2 V 2) = p (3 V 3) being a consequence of these two. Corre
spondingly, of the three ClapeyronClausius equations following from
these three relations
2V
only two are independent, while the third is a consequence of them.
If we plot the pressure of the system against its temperature,
the three equations (7.17) will represent three curves OA, OB, OC
(Fig. 13). All points of the p, Tplane which
do not lie on one of these curves correspond to a
A / singlephase state of the system. The points on
I I .*& ^n the curves OA, OB, OC represent the equilib
rium of two phases, and the point of intersection
O is the triple point. As an illustration of the
(p, r)diagram, we have taken the most com
mon case when the three phases in question are
vapor, liquid, and crystalline solid. The branch
OA representing equilibrium of the liquid and
gaseous phases ends abruptly at the critical
point A, since there is no difference between
these two states at higher temperatures. 1 Suppose that we have
vapor and liquid in equilibrium and slowly change the boiling pressure
and temperature so as to move the representative point of the system
along the curve of boiling AO towards the triple point 0. When the
point is reached, a contingency may arise similar to that discussed
in section 43 with respect to the expansion of a liquid. Under special
conditions of great purity of the material and of very careful manipu
lation, the solid phase may fail to appear and the vaporliquid combina
tion may get into a state of supercooled and unstable equilibrium
represented by the dotted continuation of the curve A 0. The slightest
shock may cause then a sudden crystallization.
1 With respect to attempts of extending the branch OA beyond the critical point,
compare: Trautz and Ader, Phys. Zs. 35, p. 446, 711, 1934; Eucken, ibidem,
p. 708.
FIG. 13. Triple point
and equilibrium of solid,
liquid, and vapor.
VII 48 PHASES OF A CHEMICALLY PURE SUBSTANCE
127
The triple point of most substances lies at pressures which are low
compared with the normal atmospheric. When they are heated at
constant atmospheric pressure, their state changes along a vertical
straight line drawn to the right of the triple point. The crystal is first
converted at a certain temperature (of fusion) into a liquid ; at a higher
temperature the liquid is vaporized. It is different, when the pressure
of the triple point is high or when the substance is heated at a low
pressure: then, the vertical lies to the left of the point and the
crystal is directly sublimated into vapor without passing through the
liquid state.
The bestknown example of the coexistence of three phases is that
of steam, water, and ice at p = 4.579 mm Hg and / = 0.0075 C.
However, researches by Tammann l and Bridgman 2 have disclosed
that, in addition to the ordinary ice (I), there exist at least four other
crystalline states of water (ices II, III, V, VI) which are all denser
than water and observable only at very high pressures. (The existence
of a modification, formerly labeled ice IV, is considered as doubtful).
The table of the several triple points, as far as they have been observed,
is as follows
TABLE 13
PHASES
t
P
FIG. 14
Ice I, liquid, vapor. . . .
0.0075 C
4.579 mm Hg
Ice I, liquid, Ice III. . .
22
2115 kg/cm
C
Ice III, liquid, Ice V. . .
17
3530 ' '
D
Ice V, liquid, Ice VI ...
f OM6
6380 ' '
E
Ice I, Ice II, Ice III...
37. 7
2170 "
F
Ice II, Ice III, IceV...
24. 3
3510 "
G
A graphical representation of the measurements carried out with
water in its different phases is given in Fig. 14.
An interesting situation was disclosed by Tammann 's investigation
of the equilibrium curves of phosphonium chloride, PH4C1 (Fig. 15).
The curve of fusion could be traced to temperatures considerably
higher than the critical point of the liquidvapor equilibrium. It seems
paradoxical that the crystalline phase can be observed at temperatures
at which the liquid does not exist or, rather, cannot be distinguished
from the vapor. A similar behavior is shown by carbon dioxide (62)
and by helium (He). The curve of fusion of CO2 was observed by
1 G. Tammann, Ann. d. Phys. 2, p. 422. 1900.
* Bridgman, Proc. Amer. Acad. 47, p. 441, 1912.
128
TEXTBOOK OF THERMODYNAMICS
VII 49
Bridgman up to / = 93.5 C and p = 12 000 kg/cm 2 , while the critical
data are t c = 31.4C, p c = 72.9 atm. The case of helium is even
more striking: Simon, Ruhemann, and Edwards 1 observed its curve
12000
10000
8000
4000
2000
^040302010 10
rc
FIG. 14. Phase equilibrium in water.
80
of fusion up to T = 42 abs and 5600 atm, i.e. solid helium can be
prepared at a temperature fully eight times as high as the critical
point of liquid helium (t c = 5.2
abs, pc = 2.25 atm). The question
is still open whether the curve of
fusion extends indefinitely to still
higher temperatures or is limited,
either by a solidliquid critical
point or in some other way. Its
experimental investigation is very
difficult as it requires the applica
10 20 30 40 50 tion of ex tremely high pressures.
100H
80
60
40
20
FIG. 15. Phase equilibrium in phos
phonium chloride.
49. Phase equilibrium of higher
order. 2 The usual case of two
phases in contact treated in section
1 may be called equilibrium of the
first order. Influenced by observations on liquid helium, Ehrenfest 3
discussed another possible case which he called equilibrium of the second
1 Simon, Ruhemann, and Edwards, Zs. phys. Chem. (B) 2, p. 340, 1929; 6,
p. 62, 1930.
2 This section may be skipped without loss of continuity.
3 P. Ehrenfest, Communications Leiden, Suppl. 756, 1933.
VII 49 PHASES OF A CHEMICALLY PURE SUBSTANCE
129
order. It may happen that the transformation takes place without
development of latent heat and without change of volume. Supposing
the molecule to be the same in the two phases,
X2 Xi = 0, AF = V2 v\ = 0.
(7.18)
This means that the two phases have the same molal heat function
and the same molal volume. Because of the fundamental relation of
equilibrium,
A$ = 0, 2  <PI = 0, (7.19)
snce
they have then also the same entropy (AS = $2 si == 0),
A5 = Qp/T, according to (4.11). Yet the phases are different because
of their other properties (for instance, the molal heats) changing dis
30
25
20
C 15
10
5
2.0 2.2 2.4 2.6 2.8 3.0K
*~T
FIG. 16. Specific heat of helium.
continuously in the transformation. This seems to happen with
liquid helium at the temperature of 2.2 K: the discontinuity in its
molal heat is given in Fig. 16 as measured in Leiden l while changes
of entropy and volume were not observed. Similar conditions are
known to obtain in many other transformations: solid methane at
20.4 K, solid oxygen at 44 K, HBr at 87 K, ammonium chloride
and bromide, etc. 2 The accuracy of these observations was in no case
sufficient for a conclusive proof that the conditions (7.18) were strictly
satisfied, but it is interesting to follow up the theoretical implications
of such a contingency.
1 W. H. Keesom and A. P. Keesom, Comm. Leiden 222 d; W. H. Keesom and
K. Clusius, 219 e; K. Clusius and A. Perlick, Zs. phys. Chemie 24, p. 313, 1934.
2 Compare reviews by M. Von Laue and A. Eucken, Phys. Zs. 35, pp. 945, 954,
1934.
130 TEXTBOOK OF THERMODYNAMICS VII 49
Taking account of (6.51) and (6.52), the conditions (7.18) can be
also written in the form
0, A*,0, (7.20)
which converts the ClausiusClapeyron equation (7.04) into an iden
tity. In order to obtain a relation between the differentials of tempera
ture and pressure, Ehrenfest proceeds to differentiate these two
equations.
where the partial differentiation with respect to T or p is indicated by
writing these variables as subscripts. He further postulates that these
two relations be compatible. It should be remembered that, according
to (5.37), (4.32)
~~ C PI
(7.22)
The elimination of dp/dt from eqs. (7.21) gives Ehrenfest's relation
for equilibrium of the second order

Serious objections against these conclusions were raised by Keesom 1
and by Von Laue. 2 They can be brought out most clearly by a geo
metrical interpretation of eqs. (7.19) and (7.20). If we regard T, p, <p
as cartesian coordinates of a threedimensional space, the equations
<p <pi(p, T) and <p = <f>z(p, T) represent two surfaces in this space.
In the ordinary first order equilibrium the condition (7.19) is satisfied
along the line of intersection of these two surfaces, as represented
graphically in Fig 17a, which gives the trace upon the (<p, T)plane.
The relation between the transformation values of p and T, given
by the ClausiusClapeyron equation (and represented graphically in
the figures of section 48), is then the projection of this line of inter
section onto the plane p, T. The additional conditions (7.20) for the
second order transformation mean that the two surfaces are in contact
1 W. H. and A. P. Keesom, Physica 1, p. 161, 1933.
2 Footnote on p. 129.
VII 49 PHASES OF A CHEMICALLY PURE SUBSTANCE 131
(Fig. 176). Von Laue points out that in this case no transformation
at all is possible. In fact, the theory of equilibrium requires A4> g
(section 36) and, therefore, the states of the system represented by the
upper branches of the curves are unstable. It will be found in the
more stable state (1) both above and below the temperature of the
contact. Von Laue suggests, therefore, that the phases in question
may have an equilibrium of the third order, i.e. the surfaces <p = <p\
(a) (b) (0
^
Xv fi
*i
FIG. 17. Equilibrium of first, second, and third order.
and <p w have contact with penetration (Fig. 17c). The mathe
matical conditions for this are
A^JT = A$ pr = A$ pp = 0, (7.24)
in addition to (7.18) and (7.19). They represent so much of a restric
tion upon the parameters of the two phases that Von Laue's suggestion
was not favorably received by the other workers in this field.
We think, however, that the topic should not be dismissed without
a fuller investigation of the possibilities of the second order equilibrium.
Ehrenfest and Von Laue make an assumption which, in our opinion,
goes too far. They differentiate eqs. (7.20), implying that the contact
exists along the whole line of intersection of the surfaces <pi, w or, at
least, along a finite part of it. There is nothing in the experimental
observations to justify this assumption. Theoretically it is also a
rather remote case, because it cannot occur, unless the properties of
the two phases happen to satisfy Ehrenfest's eq. (7.23) or Von Laue's
even more restrictive conditions (7.24). It is a far more common
occurrence that the two surfaces are in contact in a singular point, i.e.
eqs. (7.20) are satisfied just in one point whose coordinates we shall
denote by po, TQ. What is then the value of A< in a neighboring point,
say, PQ + dp, TQ + dT? Expanding with respect to dp, dT to terms
of the second order
dT, po + dp)
. . . (7.25)
132
TEXTBOOK OF THERMODYNAMICS
VII 49
Because of the conditions (7.19) and (7.20), equilibrium in the
vicinity of p, T is then determined by the condition
0.
The expression D of formula (7.23) is the discriminant of this
equation. When D < 0, the equation has no real solutions so that a
change of phase is impossible. The limiting case, D = 0, is that of
Ehrenfest's and must be excluded because of the reasons advanced
by Von Laue. But when D > 0, the equation has two real solutions.
Because of (7.22),
dT
dp
r/9AF\
lA QT J p
TQ
(7.26)
Consequently, transformations are possible, if not in the point
po, To itself, in its immediate vicinity as is illustrated in the (p, T)
diagram of Fig. 18. There are two lines in which the phases (1) and (2)
(2)
(a)
(b)
A'
(2)
"
\] (1)
FIG. 18. Vicinity of a point where
equilibrium is of second order.
FIG. 19. Special cases of first order
equilibrium.
can coexist, and they intersect in the singular point po, TQ. When
the substance is cooled at a pressure exactly coinciding with that of
the intersection (dotted line AB), no transformation takes place. If
it is cooled at a slightly different pressure (line A'B') it is transformed
first from the phase (2) into the phase (1) and, at a lower temperature,
back again into (2). Some of the experimentally measured anomalies
present features compatible with this picture: as the temperature is
changed, the specific heat jumps discontinuously and, at a slightly
lower temperature, jumps back to the initial value (Fig. 16). How
ever, it would require a more detailed experimental investigation to
establish the existence of this type of transformation beyond doubt
in view of the following considerations.
VII 49 PHASES OF A CHEMICALLY PURE SUBSTANCE 133
More important than the direct application of the preceding
results is the light they throw, indirectly, on transformations of the
first order with small but finite latent heats and volume changes.
Let po, To be a point in which A$r and A$ p are very small but finite.
In the infinitesimal vicinity of the point po, To there will hold, of
course, the ClausiusClapeyron equation. But if we consider, instead
of the differentials dp, dT t small finite increments A/>, AT, the second
order terms in (7.25) may no longer be neglected and the equation
becomes
= 0.
This represents an ellipse or a hyperbola with its vertex near the
point po, To, according to whether the discriminant (D) is negative or
positive (Fig. 19). The dotted parts of the curves have no reality
because the approximation breaks down when A/?, AT become appreci
able since then terms of third and higher order must be taken into
account. The gist of the matter is that these />, jTlines are strongly
curved. In cooling the system along the lines A'B' we meet with
conditions quite similar to those of Fig. 18. In fact, the transition
from the first to the second order of equilibrium is a continuous one
and the two cases cannot be distinguished without a complete experi
mental investigation of the whole region around the point po, TQ.
Some of the cases enumerated in the beginning of this section are,
probably, examples of this kind of equilibrium of the first order. 1
Another interesting illustration will be presented in section 137.
1 The same conclusions were drawn by A. Eucken (footnote on p. 129) from the
study of the experimental material.
CHAPTER VIII
MIXTURES OF PERFECT GASES
50. The thermodynamic potential of a gas mixture. An example
in which the general theory of equilibrium can be completely carried
through in detail is the mixture of perfect gases. This is largely due
to the fact emphasized in section 4 that the pressure of such a mixture
is equal to the sum of the partial pressures of the individual gases
p = pi+ P2 + . + p., (8.01)
where pi, . . . p 9 are computed for each gas as if it were filling the
available volume alone and the other gases were not present. Suppose
that we have in a vessel of the volume V several gaseous constituents
in the respective amounts of N\, . . . N 9 mols. Let their partial
pressures be expressed according to eq. (1.12)
Vh being the molal volume of the gas fe. In order to apply to this system
the theory of Chapter VI, we must find the thermodynamic potential
$ = JJ TS + pV, where T and p are the temperature and the total
pressure of the mixture. The physical inference from the formula
(8.01) is that there is no interaction between the gases, so that the
thermodynamic potential of each of them must be computed, as if it
were alone in the vessel, and all these potentials must be added to
obtain
9
*=y%*. (8  3)
trr
A simple and convincing way of seeing this is opened by an ideal
experiment devised by Gibbs. We give the vessel containing the
system the following construction (Fig. 20). The cylinder C fits tightly
into the cylinder C 1 and can slide within it without friction. The outer
ends of both cylinders are closed by tops AB and A'B' impenetrable for
the gases and adiabatic like the side surfaces. They have also bottoms
134
VIII 50 MIXTURES OF PERFECT GASES 135
MN and M'N' which are kept by a suitable arrangement at constant
distances / from the tops while the cylinders are sliding. However,
these bottoms are semipermeable membranes: The membrane M'N'
lets through freely the component h but is impenetrable for all the
others. On the other hand, M'N' is impenetrable only for h, offering
no resistance to the rest of the components. The volume of each
cylinder (ABMN and A'B'N'M') is the same and equal to V. We
start our considerations from the state when the cylinder ABMN is
pushed completely in, so that MN touches A'B' and the gas mixture
fills the joint volume V. We begin now to pull out this cylinder in a
reversible (infinitely slow) way. Three part volumes of the system
must now be distinguished: (1) the section above M'N' is filled with
the pure h component because this membrane is not permeable to the
other gases. (2) The middle part between M'N' and MN contains all
components. (3) The section below MN is free
of the component h because this gas is pushed up
by the membrane MN which is impermeable to it.
The gas h is, therefore, present only in sections (1)
and (2) (or in the upper cylinder). We have
emphasized that its partial pressure p\ is inde
pendent of the presence of other gases: It will
be, therefore, the same in section (1) where it is
alone and in (2) where it is one of the constitu
ents of the mixture. Only the component h A'
exercises pressure upon the top AB and the FIG. 20. Device for
bottom MN of the upper cylinder. The rest of unmixing gases,
the gases do not produce any pressure effects
on the two surfaces: on AB, because no other component is
present in the part volume (1) adjacent to it; on MN, because this
membrane is completely permeable to the other gases. We see,
therefore, that the forces of pressure on top and bottom are oppositely
equal so that the resultant force on the whole cylinder vanishes. The
process of reversibly pulling out the cylinder does not involve any
work (W = 0), provided there are no outer forces upon the system
(e.g. airfree outer space). Since the outer envelope of the system is
adiabatic (Q = 0), the reversible process goes on without change of
entropy (section 23). Moreover, it proceeds at constant internal
energy and temperature. In fact, in the general eq. (3.05) of the first
law, Q = t/2 Ui + W, the terms Qand W vanish leaving C/2 = E/i,
and in the case of perfect gases, this is equivalent to T% Ti, accord
ing to eq. (3.18) or (3.19).
Continuing the process until the membranes MN and M'N' touchy
136 TEXTBOOK OF THERMODYNAMICS VIII 50
so that the middle section (2) disappears, one can entirely separate the
component h from the mixture. In a similar way, the other compo
nents can be separated, one by one, so that, in the end, the system is
completely unmixed and each component occupies, separately, a
volume equal to V. The whole process is carried out without change
of U and S. It follows from this that the internal energy and the
entropy of a gas mixture are additive U = 2C/A, S = SS*, provided
the entropies Sh of the individual perfect gases are computed as if
each of them occupied the volume V alone. Therefore, the first two
terms of the thermodynamic potential $ = U TS + pV are addi
tive, while the additivity of the third follows from the relation (8.01),
thus completing the proof of eq. (8.03).
If we use eq. (4.20) for the molal entropy of a perfect gas, the
entropy Sh of the component h has the expression
S h  N k j<ir  R log p h
and the total entropy of the mixture becomes
5  2] N "  R log pk + s * (8  04)
Sometimes it is preferable to express 5 in terms of the total pressure
p rather than the partial pressures ph. In order to carry out the trans
formation, it is convenient to use the mol fractions defined in section 39
x h = Nk/N, (8.05)
where N = Ni + . . . + N 9 represents the total number of mols (of
the different constituents) in the mixture. In the case of perfect gases,
equations p h V = N h RT and pV = (N\ + . . . + NJRT hold, so that
ph/p = x h or
ph = x h p. (8.06)
Substituting this into (8.04)
. (8.07)
The first three terms of this expression are linear in the mol numbers
while the last shows a more complicated dependence upon them.
VIII 51 MIXTURES OF PERFECT GASES 137
In a similar way, there follows from eq. (5.41) for the molal thermo
dynamic potential
7k = RT log p h + a*(TO, (8.08)
and
, N h [RT log p k + (DL (8.09)
T^T
where
/T s* c
c ph dT  T I ~ dT + u Qh  7*0*. (8.10)
J l
or an alternative form after the substitution of (8.06).
We denote the molal potential (8.08) by the symbol ^ instead of
<ph for reasons which will become clear in the next chapter. This is
entirely permissible because, in this case, the partial thermodynamic
potentials of the mixture are identical with the molal of the pure com
ponents, as a consequence of the additivity expressed in eqs. (8.03)
and (8.09). In fact, as determined by eq. (8.08), <pn satisfies the
definition (5.38) of the partial molal thermodynamic potential.
Exercise 76. Derive the expressions of the heat function X and the work
function ^ for mixtures of gases.
Exercise 77. Check the statement that the derivative (3*/3^)p f r is identical
with (8.08). In the differentiation the total pressure is to be kept constant.
51. The mass law. If the gases of the mixture can react chemi
cally with one another, according to the formula (6.46), S^G* = 0,
the condition of equilibrium is given by eq. (6.49) or 2*>^ = 0.
Substituting for <ph the expressions (8.08), we obtain
log p h = 
We recall that the coefficients j>& of the components consumed in
the reaction have the negative sign, of those produced in it the positive.
Therefore, the physical meaning of the expressions on the right side,
when (8.10) is substituted, is as follows:
= AC P (8.12)
represents the difference of the heat capacities (at constant pressure)
of the gases produced and consumed in the reaction, or the total
change of the heat capacity due to it. According to eq. (3.27),
138 TEXTBOOK OF THERMODYNAMICS VIII 51
c p hdT + UOH = *h is the heat function of the component ft. For
this reason
Q P (8.13)
represents the total change of the heat function of the mixture AX
which is equal to the heat of reaction Q p , as we know from eq. (3.23).
The last term depending on the entropy constants of the constituents
of the mixture, we shall denote by
= R log L (8.14)
We further introduce for the whole right side of eq. (8.11) the
notation RT log K pt so that
!<* Kp = / ^ dT  Jfr + ^g L (8.15)
Within the range of temperatures, where the molal heats c p h can
be considered as constant, this becomes
log K P = ~2 log T  QL + i og /. (8 . 16)
Strictly speaking, / is well defined only in this case, unless the
lower limit of the integral in (8.15) is specified.
The condition of equilibrium (8.10) takes, therefore, the form
KW .../' = p. (8.17)
An equivalent form is obtained by replacing the partial pressures by
the molal fractions by means of eq. (8.06), ph = Xhp 9
xi*xf . */* = K, (8.18)
where K = K p /p" or
log K = log /  j, + dT  v log p t (8.19)
with the abbreviation
*=Vl + 9* + ... + 9r (8.20)
The relation (8.17) between the partial pressures of the constituent
gases or the equivalent eq. (8.18) between the molal fractions is known
as the mass law. Whatever the constitution of the mixture may be,
VIII 51 MIXTURES OF PERFECT GASES 139
at a definite temperature and pressure, the function pi*p* ... (or
Xi n x 2 n . . . ) is constant. The quantity K p (respectively, K) is,
therefore, called the equilibrium constant of the reaction.
The mass law was first deduced by the Norwegian chemists
Guldberg and Waage (1867) from statistical considerations which it
will be instructive to reproduce here. Let us write the equation (6.46)
of the reaction in the form
vid + v 2 G 2 + . . . <=* v'iG'i + v' 2 G' 2 + . . . ,
where the symbols, on one side, refer to the substances entering the
reaction and, on the other, to those resulting from it. In order that
vi, i> 2 , . . . molecules of the substances GI, G 2 , . . . enter the reaction,
they must all meet, i.e. be simultaneously present in a small volume
TO. What is the composite probability of such an event? The prob
ability of one molecule of the kind GI being in the volume TO is pro
portional to the total number of such molecules available, which, in
its turn, is proportional to the mol number N\. The probability of vi
molecules being there simultaneously is proportional to N\ l (accord
ing to the rules expounded in section 30), provided that the molecules
are statistically independent. This statistical independence implies
that there are no forces of interaction of any kind between them, and
this is true only for perfect gases. Since we have to restrict ourselves
to this case, we may make use of the property of perfect gases that
the mol number is proportional to the partial pressure and conclude
that the above probability is proportional to pi* 1 . In a similar way,
we find for the probability of v 2 molecules of the second component
being in the volume T O the expression p<P> and so on. The probability
of the composite event is, therefore,
where the constant of proportionality a may depend upon the prop
erties of the system, especially its temperature and pressure. The
reaction may, however, go also in the opposite direction producing
Pi t v 2 , . . . molecules of the substances GI, G 2t ... It is necessary for
this that i/i, 1/2, molecules of the substances G'i, G' 2 , . . . enter
the reaction and meet in a limited volume. The probability for this
is, obviously,
P' = a'fi'yS* . . .
In the state of equilibrium the number of molecules of every kind
140 TEXTBOOK OF THERMODYNAMICS VIII 52
is stationary and the process must go as often one way as the other.
The probabilities P and P' must be equal
(K p = a'/a) an equation which is identical with (8.17).
The thermodynamical derivation given above is due to Gibbs (1871)
and leads a good way beyond the result of Guldberg and Waage inas
much as it contains in eqs. (8.15) and (8.19) the explicit dependence of
the equilibrium constant upon p and T. It is sufficient to make one
measurement at a definite pressure and temperature in order to
determine the constant /. Knowing /, one can predict the equilibrium
conditions for all possible states of the mixture.
Exercise 78. Find the form of the mass law (8.17) for the following reactions:
N 2 + 3H 2  2NH, = 0; 2HBr + I,  2HI  Br a = 0; 2O  O 2  0;
2O 8  3O 2  0.
Exercise 79. Show that, in view of eqs. (4.36") and (3.26), the expression (8.15)
can be reduced to
(8.21)
J JKJ. "
which is consistent with (6.53).
52. Examples. Degree of dissociation. We have obtained two
forms of the mass law in eqs. (8.17) and (8.18). The first form is
simpler inasmuch as the quantity K p is a function of the temperature
only. All the quantities on which eq. (8.17) depends are completely
conserved if we add neutral gases to the mixture, because the partial
pressures pi, . p 9 , as well as K p , remain unaffected by this. The
second form has the advantage that it can be more readily generalized
to include systems which are not perfect gas mixtures.
The most interesting question is, usually, as to the fraction of the
original substance or substances transformed in the reaction. Let us
discuss, for instance, the oldest reaction to which this theory was
applied by Gibbs, the dissociation of nitrogen tetroxide
2NO 2  N 2 O 4 = 0. (8.22)
The coefficients v are in this case v Nl04 1> ^NO* 2, v = 1.
Suppose that No moles of N 2 C>4 are brought into a vessel at a very
low temperature when it is quite undissociated : what mol numbers
shall we observe after heating the gas to the temperature T at the
pressure p? If we denote by { the fraction of the N 2 C>4 molecules
which are dissociated (degree of dissociation) the mol number of this
VIII 52 MIXTURES OF PERFECT GASES 141
component is ^ N2 o 4 = ^o(l ) To every dissociated molecule
there appear in the mixture two molecules of NC>2. The mol number
of nitrogen dioxide is, therefore, jV N o, = 2No, and the sum
N = # Nl04 + WNO, = N (l + ) Recalling the definition (8.05) of
the mol fractions and substituting into (8.18), we find
K = K p /p. (8.23)
I? 2
We see from this that the degree of dissociation is small when K is
small, i.e. according to (8.19), at low temperatures or high pressures.
Under the opposite conditions of high temperature or low pressure,
approaches unity.
The degree of dissociation being known, it is easy to compute the
density, which is also accessible to direct measurement. The density
of the mixture A, obviously, stands to the density 5 of the undissoci
ated component N2O4 in the inverse ratio of the numbers of molecules:
A/6 = No/N, whence
A =
Accurate measurements on the dissociation of nitrogen tetroxide
are due to Bodenstein. 1 This gas can be regarded as approximately
perfect only when the pressures are not too high and the temperatures
not too low. From observations under these conditions Bodenstein
derives the formula
logic K p =  ~ + 1.75 logio T + 0.00483r
 7.144 X 10 6 r 2 + 5.943, (8.25)
when the pressure is expressed in mm Hg. The agreement with obser
vations can be judged from Table 14.
TABLE 14
CHEMICAL EQUILIBRIUM IN N 2 O 4
T(abs) 282. 5 285. 7 289. 5 293. 305. 9 323. 9
logio K p (obs.) 1.487 1.608 1.745 1.859 2.294 2.838
logio Kp(calc.).... 1.475 1.600 1.741 1.889 2.286 2.838
T(abs) 334. 8 342. 7 352. 5 361. 9 373. 9 387. 4
logio K p (obs.) 3.135 3.326 3.577 3.784 4.027 4.293
logio K p (calc.). ... 3.135 3.341 3.577 3.794 4.032 4.323
1 M. Bodenstein, Zs. phys. Chemie 100, p. 74, 1922.
142
TEXTBOOK OF THERMODYNAMICS
VIII 52
Bodenstein's formula represents also fairly well the older measure
ments by Natanson. 1 Although not very accurate, they offer us exam
ples of how the degree of dissociation varies with the pressure at
constant temperature. One of the series was measured at / = 49. 7 C,
for which temperature (8.25) gives K p = 643. The normal density
of N 2 O4 is d = 3.180, whence the actual density (reduced to normal)
is calculated by means of eq. (8.24).
TABLE 15
DISSOCIATION OF N 2 O 4
Pmm
Kp/tp
*
A (calc.)
A (obs.)
1.000
1.590
26.80
5.997
0.926
1.651
1.663
93.75
1.715
0.795
1.772
1.788
182.69
0.880
0.684
1.897
1.894
261.37
0.615
0.615
1.969
1.963
497.75
0.323
0.494
2.129
2.144
A reaction of great practical interest is the dissociation of water
vapor
2H 2 + O 2  2H 2 O = 0,
with the coefficients V H , = 2, v^ = 1, v Ht0 = 2, v = 1. We suppose
again that we put into the reaction vessel JVo undissociated water
molecules (at a low temperature). At a higher temperature dissoci
ation ensues and the mol numbers of the three components become
Equation (8.18) gives, then,
1
*>
P'
(8.26)
2 (i  )2 (1 + f
According to Siegel 2 the heat of this reaction is
Q 9  113 820 + 2.6Sr  4.41 X 10~ 4 r 2 + 1.252 X 10~ 6 r 3
9.12 X 10 10 r 4 + 4.36 X lO^r 6 , (8.27)
1 E. and L. Natanson, Ann. Physik u. Chemie 24, p. 454, 1885.
*W. Siegel, Zs. phys. Chemie 87, p. 641, 1914. The value of the constant
1.08 is due to Nernst and Wartenberg.
VIII 53
MIXTURES OP PERFECT GASES
143
and the equilibrium constant
24900
logic K P = ~
+ 1.335 logio r  0.965 X 10~ 4 r
+ 0.137 X 10~ 6 r 2  0.665 X 10 10 r 3 + 0.191 X 10 17 r 6  1.08. (8.28)
Nernst 1 gives the following comparison of observed data of the
degree of ionization with those calculated from this formula (at
atmospheric pressure) :
TABLE 16
DISSOCIATION OF WATER VAPOR
T (abs)
100$ (calc.)
100 (obs.)
T (abs)
lOOt (calc.)
100$ (obs.)
290
4.66X10 2 '
4. 6  4. 8 X 10~ 2
1705
0.107
0.102
700
5.4 X10
7.6X10*
2155
1.18
1.18
1300
0.0029
0.0027
2257
1.76
1.77
1397
0.0085
0.0078
2337
2.7
3.8
1480
0.0186
0.0189
2507
4.1
4.5
1500
0.0225
0.0197
2684
6.6
6.2
1561
0.0369
0.034
2731
7.4
8.2
3092
15.4
13.0
The agreement is remarkable in view of the width of the range and
of the fact that the observations were obtained by many different
observers and methods.
Exercise 80. Find the equation for the degree of dissociation, analogous to
(8.23) and (8.26), for the reactions: 3H 2 + N 2 = 2NH,, Hg + I 2  HgI 2 .
Exercise 81. Calculate for hydrogen iodide from eq. (8.30) at the pressure
p = 100 mm and at the temperatures of Table 17.
53. Influence of the excess of one component. Another inter
esting reaction is the dissociation of hydrogen iodide
H 2 + I 2  2HI  0,
?H, = 1, "i, = 1, v
 2, v = 0.
(8.29)
The equation v = means that the number of molecules is not
changed in the reaction. The equilibrium constant (8.19), then, does
1 W. Nernst, Theoretische Chemie, p. 775. llth edition. Stuttgart 1926.
144 TEXTBOOK OF THERMODYNAMICS VIII 53
not contain p, so that the molal fractions and the degree of dissociation
are entirely independent of the total pressure. Suppose we bring
into a vessel (at a low temperature) N\ mols of HI and N% mols of Ha
and heat the mixture to the temperature T. This assumption will
permit us to study the influence of the excess of one component (Ha)
upon the degree of dissociation . As in the preceding case,
N m = Ni(l ). Each dissociated molecule of HI produces half
a molecule of Ha and la, respectively: therefore, JV Ia = %Ni and
^Ha = 2^1 { + A/a, while the sum, N = NI + NZ, is independent
of the degree of dissociation. The substitution into (8.17) gives
+
* *' (8 ' 30)
When the temperature is low, K is small and, with it, the degree of
dissociation . We can take then for the denominator the approximate
value 1 and see that the numerical value of greatly depends upon
the term 2Nz/N{. When there is no excess hydrogen (N% = 0), we
find = 2K*\ on the other hand, when 2Nz/Ni is not small, we can
neglect , compared with it, obtaining = 2KN\/N%, a number
which is much smaller: The excess of one of the products of dissociation
seriously depresses its degree. When K is very large (high temper
atures), approaches unity. As an approximation, we can substitute
= 1 in the numerator with the result 1 = (1 + 2N 2 /Ni)*/2K*.
Qualitatively the presence of the term 2Nz/N\ has still a depressing
influence. However, quantitatively sinks appreciably below 1
only when 2N%/N\ begins to approach K in its order of mag
nitude.
The physical reason of this depressive action is, of course, that the
excess of Ha gives to the iodine, produced in the dissociation process,
a larger chance of reacting with hydrogen and of being transformed
back into hydrogen iodide. This simple explanation makes it clear
that we have here a general phenomenon which occurs, under similar
conditions, in all cases of dissociation.
The equilibrium constant of the reaction (8.29) is well represented
by the formula 1
540 4
logic K = logio K p =   + O.S03 logio T  2.350. (8.31)
1 Nernst, Zs. Electrochcmic 18, p. 687, 1909; K. Vogcl v. Falkenstein, Zs. phys.
Chcmie72, p. 113, 1910.
VIII 54 MIXTURES OF PERFECT GASES
Hence and from (8.30) we obtain Table 17.
TABLE 17
DISSOCIATION OF HI IN THE PRESENCE OF AN EXCESS OF H s
145
Nt/Ni =
Nt/Ni  1
r(abs)
logio K
100$ (calc.)
100$ (obs.)
100$ (calc.)
600
1.856
19.1
18.9
2.6
700
1.691
22.2
21.5
3.7
800
1.565
24.9
24.7
4.8
1295
 1 . 202
33.3
32.9
9.8
1490
1.117
35.6
37.5
11.3
The depressive action is very marked, indeed.
Exercise 82. Calculate from (8.31) with the help of (8.21) the heat of the
reaction (8.29) at 1000 K. Do the same for the reaction (8.22) by means of the
formula (8.25).
Exercise 83. Derive the generalization of the formula (8.26) when there is an
excess of H a , and estimate its depressive influence.
54. Influence of neutral gases. In the absence of neutral gases
the mol fraction of each component is given by the expression
x k = N k /(Ni + ...+NJ. (8.32)
On the other hand, if there are in the mixture other gases than
those taking part in the reaction and the sum of the mol numbers of
these neutral gases is NQ, the molal pressure of the same component
becomes
Ni+...+N.). (8.33)
Comparing these two relations we find
*>k = *k(Ni + . + N,)/(N + Ni + ...+NJ. (8.34)
The expression of the mass law in the presence of neutral gases is
in accordance with eq. (8.17)
K
**l
or substituting (8.34)
146 TEXTBOOK OF THERMODYNAMICS VIII 55
The addition of neutral gases at P = const has, therefore, the same
effect as an increase of the equilibrium constant. This influence is
opposite to that of the excess of an active component because to a
larger constant K corresponds a higher degree of dissociation. The
presence of neutral gases in general advances the degree of dissociation,
provided that the total pressure p remains the same. Of course, in the
particular case v = 0, neutral gases have no effect.
We see from the definition (8.19) of the function K that the same
result could be achieved by reducing the total pressure p of the mixture
in the proportion (Ni + . . . + N ff )/(N Q + . . . + N ff ). In a certain
sense, the influence of the neutral gases is, therefore, only apparent.
It comes from the fact that we refer the effect to constant total pressure
to which both the active and the neutral gases contribute, so that the
partial pressures of the active components are reduced in precisely
the same proportion. This corresponds to the usual experimental
conditions in which the total pressure p is the fundamental datum.
If the question were put as to the degree of ionization in a box of
constant volume, the introduction of additional neutral gases into
that box would have no influence whatever on it.
Exercise 84. What is the increase of volume of the gas mixture in a reaction
in which the sum of the mol numbers increases by v? Use the equation of perfect
gases (1.18) and check that the result is consistent with (6.51).
55. Mutual influence of two simultaneous reactions. Suppose
that we mix at a low temperature N\ mols of hydrogen iodide (HI)
and N2 mols of hydrogen bromide (HBr) and slowly heat the mixture.
What will be its composition at a temperature T? In addition to the
reaction (8.29), there can take place also the reaction of dissociation
of hydrogen bromide
H 2 + Br 2  2HBr = 0,
J
1. )
"Hj = 1 "Bri = 1 ^HBr = 2,
Accordingly there must be simultaneously satisfied two equations
of the type (8.18)
(8 37)
9
HI x HBr
We can use the expressions of section 53 for the mol numbers of
HI and of 12, in terms of the degree of dissociation 1 of hydrogen
iodide: N m = Ni(l  1), N lt = jAih Similarly we can write
(denoting the degree of dissociation of HBr by {2) : NSB T =* JVa(l {2),
, since the two reactions are precisely of the same type.
VIII 55
MIXTURES OF PERFECT GASES
147
As to the mol number of hydrogen, it comes from two sources: each
dissociated molecule of both HI and HBr produces half a molecule of
H2, whence JV H , = i(^ii + N 2 2 ). Substituting into (&.QS) and
(8.37)
4 (1 i) 2 ' 4 (1 &) 2
It is apparent from these equations that the mutual influence of
the two reactions results in the decrease of the degrees of ionization of
both HI and HBr. The reason of this is easy to understand: either
of the reactions liberates hydrogen and creates an excess of this com
ponent for the other process, exercising upon it a depressive action
(compare section S3). In general, there exists still another mutual
influence: If the total pressure is kept constant, the presence of the
second gas (and of the products of its dissociation) decreases the partial
pressure of the first and, in this way, produces the same effect of
enhancing the dissociation as a neutral gas. We have already said in
the preceding section that this effect is trivial: Therefore, we have
selected here two reactions in which it does not exist because they are
not pressure sensitive. They permit us to study the mutual depressive
influence in its pure form.
The equilibrium constant of the dissociation of HBr can be repre
sented as follows
5223
logio K 2 =   + 0.553 logic T  2.72. (8.38)
At temperatures below 2000 K, 2 is, therefore, very much smaller
than ft. If NI and N 2 are of the same order of magnitude, the disso
ciation of HI is not appreciably affected by the presence of HBr, and
the values of 1 remain the same as in Table 17. On the other hand,
the dissociation of HBr is greatly depressed, as is illustrated by
Table 18.
TABLE 18
DISSOCIATION OF HBr IN THE PRESENCE OF HI
Ni =
tfl  Nl
T (abs)
logio KI
100*2
100*,
600
9.889
0.002
1.8 X 10
1000
6.284
0.14
3.4 X 10*
1490
4.470
1.16
0.12
148 TEXTBOOK OF THERMODYNAMICS VIII 55
Exercise 85. N mols of HI are brought in a vessel and heated to a high tem
perature where the reactions (8.29) and 21  I* = must be taken into account
simultaneously. Denote the degree of dissociation of HI by 1 and that of 1 2 by 2 .
The equilibrium constant for the dissociation of I 2 is (Bodenstein, Zs. Electrochem.
22, p. 338, 1916):
7550
T
K p   ^~ + 1.75 log, T  4.09 X 10* T + 4.726 X 10~ 8 T  0.440.
Calculate gs under the above conditions and in pure It at the temperature 2000 K.
CHAPTER IX
DILUTE SOLUTIONS
56. The concept of a solution. We call a solution any homogene
ous system which consists of two or more chemically pure substances,
no matter whether it is solid, liquid, or gaseous. Solutions must be
distinguished from chemical compounds, on one hand, and (in general)
from mixtures, on the other. A chemical compound consists of mole
cules of only one kind; therefore, the masses of the constituents of
which it is formed stand in a definite ratio. From the point of view
of the phase rule (section 41) it is a system with one single independent
component or a pure substance: if the amount of one constituent is
given, those of all the others are uniquely determined by it. It is
quite different with the solution; the molecules or atoms of the con
stituents do not lose their individuality in it, and their amounts can
assume any arbitrary ratio, within certain limits. The composition
of the solutions being continuously variable, they must be regarded as
homogeneous systems with several independent components.
The difference from mixtures is most clearly brought out in the
case of solid solutions. We know, at present, that a solid is always
crystalline, and we say that two substances are mutually soluble, in
the solid state, when they form a joint crystalline lattice, or a mixed
crystal in which the lattice points are occupied, in part, by atoms
(or molecules) of the one substance, in part, of the other. The solu
bility of solids is an exception, although several pairs of metals are
known which are mutually soluble in all proportions, e.g. Ag Au,
Ag Pd, Co Ni, Cu Ni, Cu Pd, Au Pt, Bi Sb, etc. The metals
of each pair are of a closely analogous chemical nature belonging to the
same group of the periodic system (or sometimes to two adjacent
groups). Of non elementary solids which are mutually soluble in
all proportions we could mention AgCl NaCl, PbBr2 PbCb,
SnCb PbCb, etc. More common is the case of limited solubility:
the crystalline structure of a metal is preserved if a fraction of the
lattice points (up to a certain limit) is replaced by atoms of another
metal, usually from the same or the adjacent group of the natural
system. The lattice cannot accommodate more than the limiting
fraction without breaking down. Such cases are Cu Ag, Au Ni,
149
ISO TEXTBOOK OF THERMODYNAMICS IX 56
Bi Pb, Sn Pb, etc. Of course, limited mutual solubility occurs
also with pairs of solids of a nonmetallic nature: KNOa NaNOa,
KNOa TiNOa, etc. All other cases, viz. those in which the con
stituents of an alloy do not form a joint crystal, must be classed as
mixtures. For instance, the microscopical examination of Ag Pb
alloys shows that they are agglomerations of tiny crystals of pure silver
and pure lead. Similar conditions prevail with most other alloys
(e.g. Bi Cd, Pb Sb, also KC1 AgCl, ice AgCl, etc.). We have
seen in section 39 that, as far as the thermodynamical applications and
conclusions are concerned, the size of the phases is immaterial; there
fore, an alloy or other mixture of this sort must be regarded as a sys
tem of two or more coexisting phases, even if these phases are of
microscopical dimensions. Unlike a mixture, a true solution is homo
geneous and represents a single phase.
The conditions in a liquid solution are similar. A liquid can, in
general, dissolve other liquids, solids, or gases, the true solution being
homogeneous and characterized by a uniform distribution over its
whole volume of the molecules or atoms of all the constituents. In
the case of liquids mutual solubility is not an exception but a common
phenomenon. At the same time there are numerous pairs of liquids
which exhibit in certain ranges of temperature a complete (e.g. water
carbon tetrachloride) or a limited (water ether, water phenol, etc.)
mutual insolubility. The true or molecular solutions must be distin
guished from the mixtures which in the case of liquids are called sus
pensions, emulsions, and colloidal solutions. Like the solid mixtures
they are inhomogeneous and represent two or more phases intermixed
in a finer or coarser degree of dispersion.
In the case of gases, the difference between a solution and a mix
ture disappears, since all gases are mutually soluble or miscible, with
out restriction, forming a homogeneous system in which the molecules
of all constituents are uniformly distributed.
The process of forming a solution is accompanied by mechanical
and thermal effects which show that there is a certain degree of inter
action between the molecules of the different constituents. When we
dissolve ethyl alcohol and water in equal parts, there is a considerable
contraction, the volume of the solution being about 5% smaller than
the sum of the volumes of the free components (compare section 32).
At the same time, a certain amount of " heat of solution " is devel
oped, because the internal energy of the solution is not equal to the
sum of the internal energies of the alcohol and water forming it.
Both effects are characteristic of the general case although their
numerical values vary widely, so that the heat of solution may be
1X57 DILUTE SOLUTIONS 151
either positive or negative. There are certain pairs of liquids known
(e.g. benzenetoluene) which form solutions with very little diminution
of volume and development of heat. This usually occurs in the case
of a close chemical resemblance of the constituents so that the inter
action between atoms of different kinds is nearly the same as that
between atoms of the same kind. We can, therefore, imagine, as an
ideal case, a solution in which the volumes, as well as the internal
energies, are strictly additive and no change occurs in these quantities
when the solution is formed. Following G. N. Lewis 1 we shall call it
a perfect solution: although the perfect solution may not occur in
nature in its pure form, it is a useful fiction as it will help us to bring
out, theoretically, some typical properties of real solutions.
57. Dilute solutions. One of the components of a solution, usually
the most plentiful is called the solvent, the others the solutes. In this
chapter we shall discuss the case when the mol number of the solvent
JVo is much larger than those of the solutes (Ni, N 2 , . . . N ). The
thermodynamical properties of every system are determined by the
characteristic functions, especially by the thermodynamic potential
which, in turn, depends on the internal energy and the volume. A
theoretical treatment of the problem requires, therefore, some knowl
edge about the analytical structure of the functions U and V, par
ticularly with respect to their dependence upon the mol numbers Nk
of the components. We have seen in section 39 that the functions
U, V, 5, $ are homogeneous of the first degree with respect to the mol
numbers. There obtains, therefore, the equation analogous to (6.13)
, T, No,... N.) = U(p,T,
Taking for e the special value e = I/No, we find
This equation (and a similar one for V/No) is quite general and
does not contain any restrictive assumptions. However, we shall
introduce now a restriction limiting the generality of the function /,
in that we assume that it can be expanded into a convergent multiple power
series with respect to the ratios NI/NQ, . . . N ff /No*
U _Ni , .N.
 WO + TT 1 + + Tf U,
NQ NQ NQ
1 G. N. Lewis, J. Am. Chem. Soc. 30, p. 668, 1908,
152 TEXTBOOK OF THERMODYNAMICS IX 57
We do not claim that this condition is always satisfied, and we
shall treat in section 115 a notable exception to it. However, the large
majority of solutions do conform to eq. (9.02), and the present chapter
will be devoted to this case. The coefficients u h , u h i, ... of the expres
sion (9.02) are functions of p and T independent of the mol numbers.
In the analogous expansion of V/No we denote the coefficients by
Vh, Vhi, etc.
We call the solution dilute if, within the desired accuracy, the terms
of second and higher order of the expansion can be neglected so that
U and V are expressed by the linear forms
[7 = NOUO + NM + ... + N ff u ff ,
V = NOV O + Ni vi + . . . + N 9 v w .
(9.03)
These equations contain, of course, the limiting case of the pure
solvent (Ni = ... = N ff = 0), when they take the form U = NQUQ,
V = NQVQ. The physical meaning of the coefficients UQ, VQ is, there
fore, the molal internal energy and the molal volume of the pure solvent.
The mol number NO, representing the solvent, cannot be put equal to
zero: therefore, the simplest form of the equations in which N h does not
vanish is U = NQUQ + NhUh It corresponds to a solution of only
the component (ti) in the solvent. Consequently, we can say that the
coefficients UH must depend on the properties both of the solute h and
of the solvent. They may be called the molal energy and volume of
the solute in the solution, but they will be, in general, different from
the same quantities characterizing the solute in its free state. When
another mol of the solvent is added to the dilute solution (9.03), its
internal energy and volume become
U' = (No + l)o + Niui + . . . + N u ff ,
V = (No + l)0o + Nivi + . . . +
so that U' = U + wo, V = V + VQ. These are exactly the same values
which the solution and the extra mol of the solvent have before they
are brought together: there is no development of heat of solution and
no change of volume when the solution is diluted still further. This
additivity would not obtain if the second order terms of eq. (9.02) had
to be taken into consideration: it can be taken as the experimental
criterion of the diluteness of a solution.
According to the definition of perfect solutions given at the end of
the preceding section, the expressions (9.03) apply also to them.
However, these solutions need not be dilute and the distinction
between solvent and solutes disappears in them. Any of the compo
1X58 DILUTE SOLUTIONS 153
nents can be regarded as the solvent and, therefore, all the coefficients
u h , Vh are identical with the molal internal energies and volumes of the
pure components. This fact guarantees the additivity of the func
tions U and V which is the essential characteristic of a perfect solution.
58. Entropy of dilute and perfect solutions. The differential of
the entropy is defined by the second law of thermodynamics as
dS = (dU + pdV}/T. The differentials dU and dV are taken for a
definite composition of the system, so that the mol numbers N h must be
regarded as constants with respect to the differentiation. Substitut
ing the expressions (9.03), we obtain
 N h (du h + pdv h ) = N ds > ( 9  04 )
ft
if we denote, for short, ds k = (du h + pdv h )/T. With respect to s h we
must repeat what was said in the preceding section with respect to
u h , v h  The quantities s h are functions of pressure and temperature
and may be called the molal entropies of the components while in the
solution. They coincide with the molal entropies in the pure state
only for the solvent and for the components of perfect solutions, and
are different from them in all other cases.
The total entropy 5, obtained by integrating eq. (9.04), will con
tain a constant of integration C, which is constant only in so far as it is
independent of the variables p, T. However, it may depend on the
mol numbers Nh (which are not variables as far as the integration is
concerned), and it can be interpreted as the entropy of mixing the
components
a
N h s k (p, T) + C(N Q , Ni,... NJ. (9.05)
This expression is entirely independent of the state of aggregation
of the system and applies to solid and liquid solutions and to gas
mixtures. If we had a complete knowledge of the equations of state
and of the energies involved, including the continuity of the states, we
could give the explicit analytical form of the functions s h (p, T) cover
ing all cases : to go from one state of aggregation to another, it would
be sufficient to change the value of the arguments p and T, as we do it
in an inaccurate way, limited to pure substances, in the Van der Waals
theory. These considerations show us that the form of the function
C(N , . . . NJ determined for one state of the system will be valid
for all other states. We did find, however, the expression of the
entropy of a mixture of perfect gases: If we admit that all substances
154 TEXTBOOK OF THERMODYNAMICS IX 59
of nature have a finite vapor pressure, be it in some cases ever so
small, there is the theoretical possibility of vaporizing any solution
into a mixture of gases. Moreover, by reducing the pressure, these
gases can be made to obey, with any degree of approximation, the laws
of perfect gases. It follows from this argument that the expression
(8.07) of the entropy of a mixture of perfect gases is only a special case
of the formula (9.05) and must have the same analytical structure.
This is, in fact, borne out by comparing them : there are in (8.07) terms
linear in the mol numbers Nh which correspond to the sum in eq. (9.05),
and terms containing the factors log x h = log (N h /N). The totality
of the latter terms must, obviously, be identified with the function C
9
C(N , Ni, ... NJ =  RY] N h log x* (9.06)
The possibility of this identification rests, entirely, on the hypoth
esis made in the preceding section that the functions U and V can be
represented as the power series (9.02) and do not contain logarithmic
terms. This assumption about the analytical form of U and V is, for
the purposes of the present theory, equivalent to the conclusions which
we shall deduce from it in the following sections, especially to Henry's
law. Van't Hoff who originated the theory of dilute solutions based
it on the validity of Henry's law. We prefer the above formulation
(which is due to Planck) because, on the one hand, it is more general,
being capable of extension by taking into consideration higher powers
of the series (9.02). On the other hand, it paves the way to the under
standing of the reasons why in certain dilute solutions the laws deduced
in this chapter are not satisfied (section 115).
59. General conditions of equilibrium. Equations (9.03) for the
internal energy and volume, together with the entropy formula (9.05),
(9.06), are sufficient to obtain the expression of the thermodynamic
potential $ = U  TS + pV, viz.
9
* = 2 Nk(vk + RT log **> (9  07)
ftO
the quantities <ph being defined as
9h = Uh Ts h + PVK, (9.08)
they are functions of p and T, independent of the mol numbers N h .
Being built up from the functions A, s*, VA, they have the properties
which we have pointed out as pertaining to these quantities. We may
call them the molal thermodynamic potentials of the components in
1X59 DILUTE SOLUTIONS 155
the solution. However, they are identical with the potentials of the
pure substances only in the cases of the components in perfect solutions
and of the solvent in dilute ones. In the case of solutes in dilute
solutions they depend also on the properties of the solvent and are,
therefore, different from the potentials of the same substances in their
free state.
The form (9.07) of the thermodynamic potential applies not only
to dilute and perfect solutions but also to mixtures of perfect gases
and to single phases of chemically pure substances. In fact, the com
parison with the expression (8.09) shows that, in the case of perfect
gases, <ph must be defined as
<p h = <p h  RT log x h = RT log p + u h (T). (9.09)
It differs from ^ of (8.08) in that p is substituted instead of the
partial pressure ph On the other hand, a chemically pure substance,
can be regarded as the limiting case of a solution without solutes
(XQ = 1, log #o = 0). We can, therefore, apply our formulas to a
heterogeneous system consisting of any number (a) of phases in equi
librium, each having one of these four characters: dilute solution, per
fect solution, mixture of gases, pure substance. If we designate the
phase by an upper index j, the total thermodynamic potential of the
system will be
izr log **<*>]. (9.10)
T^T ft^O
The comparison with the general expression (6.14) shows that the
partial thermodynamic potentials are generally $^ = <p^ } {RT log x^**
We can apply to the last equation the general theory of equilibrium of
section 42. The most general process possible in the system is given
by eq. (6.48)
The equation of equilibrium corresponding to it is, according to
(6.50),
r Ct 9 1
A*  RT I ]T ]T v h (f) log xP  log K\ = 0, (9.12)
L j1 ft0 J
or
a ff
/ j / ^ vt^ logXh^ = log K, (9.13)
jl /0
156 TEXTBOOK OF THERMODYNAMICS 1X59
where the equilibrium constant K is an abbreviation defined by
V, (9.14)
which, in the particular case of a mixture of perfect gases, can be
reduced to the form (8.19), as we shall see below. In application to
dilute solutions the formula (9.13) is due to Van't Hoff.
The left side of eq. (9.13) does not explicitly depend on temper
ature and pressure, while K is a function of p and T only, being inde
pendent of the mol numbers. An indication as to the form of the func
tional dependence of the equilibrium constant upon p and T is con
tained in eqs. (6.51) and (6.53) which give the partial derivatives of
A$ with respect to T and p. Differentiating the expression (9.12)
partially, we find
(9.16)
RT 2>
where AF and Q p are the change of volume and the heat of reaction
in the process characterized by eq. (9.11).
These equations are often called rules of displacement of equilibrium,
inasmuch as a change in the equilibrium constant leads to a displace
ment in the mol fractions of the components, according to eq. (9.13).
The first of them expresses the following fact:
When a system is compressed at constant temperature, the equilibrium
is displaced in the direction of a decrease of volume. I.e. the less volu
minous components are enhanced at the expense of the more volu
minous.
The other equation contains the rule:
When a system is heated at constant pressure the equilibrium is
displaced in the direction of heat absorption. Le. the components of
higher heat function are enhanced at the expense of those of lower.
This rule was first pointed out by Van t'Hoff and called by him the
principle of mobile equilibrium.
Exercise 86. From the definitions (9.08), (9.04), (9.03) of the quantities
<?h, Sh, Uh, Vh check the statement that they satisfy the same differential relations as
the corresponding functions in the free state of the substances, in particular that they
obey eqs. (5.37). Having ascertained this, express A V and Q p in terms of t% and 5%
from (9.15), (9.16), and (9.14).
IX (50 DILUTE SOLUTIONS 157
60. Nernst's distribution law. Let us consider the particular
case when the process (9.11) consists in the transfer of j>& (2) mols of the
component h from the phase (2) into the phase (1), where they may
form a different number of mols v h (l \ owing to association or dissocia
tion. It will simplify matters and lessen the chances of misapprehen
sion if we define both v h (l) and ^ (2) as positive and recognize the fact
that the phase (2) undergoes a loss by writing
and
A (I) log* A < l) ^ (2) log**< 2 >=log^ = [^W^^W 2 ']/^. (9.17)
Moreover, we shall often drop the superscript (2) to simplify writ
ing. With the abbreviation,
^=^ (1) /^ (2) =M^ (2) /M. (1) , (9.18)
denoting, according to (6.42), the ratio of molecular weights, this
becomes
log [(xW/Xk] = (log^)A (2) =  [W n  * h (2) ]/RT. (9.19)
In particular, when the process takes place without change of
molecule, we have ^ (1) = Vh (2) = gh= 1 and
/** (2) ) = log K h =  fo<  ^ (2) )/#r. (9.20)
In many textbooks these equations are given only in the simpler
form (9.20). Indeed, this can be justified on the following grounds.
Even if the vapor of the component h has a different, say dissociated,
molecule, the dissociation is theoretically never complete: according
to Chapter VIII, there must exist a certain fraction of undissociated
molecules, be it ever so small. Equation (9.20) gives then the equi
librium of the solution with the undissociated portion of the vapor
and permits to determine its mol fraction, whereas the mol fraction
of the main (dissociated) part of the component h can be obtained, in
a secondary way, by using the theory of Chapter VIII. However, the
more general eq. (9.19) is preferable from the practical point of view,
as it gives directly the equilibrium with either of the two fractions.
It permits one to describe by a single formula a much wider variety of
observational results, a fact which more than compensates for its
slightly more complicated form.
Equations (9.19) and (9.20) are the fundamental relations in the
theory of equilibrium of phases of dilute and perfect solutions. A num
158 TEXTBOOK OF THERMODYNAMICS 1X60
ber of important laws can be derived from them. The second equa
tion can be written in the form
*A a) /*A (2) = K h . (9.21)
The ratio of the molal fractions of any component in two phases is a
function of temperature and pressure only, provided the molecule of the
component is the same in both phases. Suppose, for instance, that
the phases (1) and (2) mainly consist of two solvents spread one on
top of the other (as would be the case with water and carbon tetra
chloride). Let the same solute h be dissolved in both solvents: the
quotient ^ (1) /^ <2) is then called the coefficient of distribution in the two
phases. We see from eq. (9.21) that the coefficient of distribution is
independent of the concentration of the solute in either phase, a law
first pointed out by Nernst in 1891.
The experimental data are, usually, expressed not in mol fractions
but in mols of the solute per 1000 cm 3 of the solution, denoted by A^
Because of the dilution, Ah can be regarded as proportional to XH\
therefore, the ratio K' = Ah (1) /A h (2} must be also constant. As an
illustration we may take iodine (12) dissolved in carbon tetrachloride
(2) and in water (1) at 25 C. 1
TABLE 19
DISTRIBUTION RATIOES
A ........ 0.02 0.04 0.06 0.08 0.09 0.10
#'0.5 ____ 85.1 85.2 85.4 86.0 86.4 87.5
In the general case, when the molecular weight of the solute is dif
ferent in the two solvents, eq. (9.19) gives
(acny/scn. K I/V( *\ (9.22)
For instance, benzoic acid has in benzene (1) a molecular weight
twice as high as in water (j> (1) = 1, j/ (2) = 2, g = J). Another example
is trichloroacetic acid (C2H ClaC^) dissolved in water (2) and in ethyl
ether (1) at 25 C. The measured data are as in Table 20. 2
TABLE 20
A .............. 0.001 0.005 0.010 0.015
A .............. 0.0021 0.0048 0.0068 0.0083
15.0 14.7 14.7 14.7
1 Washburn and Strachan, J. Am. Chem. Soc. 35, p. 681, 1913; Linhart, ibidem,
40, p. 158, 1918.
1 Smith, J. Phys. Chemistry 25, pp. 605, 616, 1921.
1X62 DILUTE SOLUTIONS 159
It appears from this that g  or ji (1) = ZM^. The molecule is
normal in the aqueous solution but has twice the normal weight in
ether as a solvent. The results are only rarely as clean cut as this: in
many solvents the association is partial, resulting in an effective
molecular weight which is a nonintegral multiple of the normal.
Exercise 87. The following values were measured in the case of Brj dissolved
in bromoform (1) and in water (2) at t = 25 C.
4 (1) = 0.125 0.25 0.50 0.75 1.00
100 X A = 0.193 0.382 0.750 1.11 1.47.
Show that the distribution law is approximately satisfied.
61. Influence of a neutral atmosphere on the vapor pressure.
We label the condensed phase (2) and the atmosphere above it (1),
and we assume that the gases of the atmosphere are but little dissolved
in the condensed phase so that it can be considered as pure (x^ 1),
with a sufficient approximation. The fundamental eq. (9.19), applied
to the vaporization of the condensed substance, gives
log *,> = fak>/ft  n]/RT. (9.23)
The potential <p h (l> relating to the vapor has the form (9.09) and
contains the term RT log p which can be brought to the left side and
combined with x h (l \ Noticing that x h ( "p = ph (l) (the partial pressure
of the vapor in the atmosphere), we have
log p k  W/ft  k(T)]/RT t (9.24)
where w A (r) is the remaining part of ^ (1) and a function of the
temperature only.
The terms on the right depend only on the properties of the com,
ponent h and not on the rest of the atmosphere. In the case when
the condensed phase is in equilibrium with its pure vapor, the left side
goes over into log p (l) (where p (l) is the total pressure) while the right
side remains the same. We see, therefore, that the atmosphere does not
influence the vapor pressure: the partial vapor pressure in the atmos
phere is identical with the total vapor pressure in its absence.
62. Henry's law. In this section we shall apply the fundamental
formulas (9.19) and (9.20) to the equilibrium of a liquid and a gaseous
phase labeled, respectively, (2) and (1). The question we shall ask
here is about the mol fraction x h of a dissolved component (gas) in
the condensed phase if its partial pressure in the adjacent atmsophere
is given. We may say that the condensed phase (2) is here preferred
inasmuch as we are particularly interested in what occurs in it. It
160 TEXTBOOK OF THERMODYNAMICS 1X62
will be well to bring the preferred phase into prominence by dropping
the superscript (2) relating to it: we shall write, therefore, XH instead
of Xk (2 \ The process to be considered is the transfer of v h mols of the
component h from the liquid solution into the gaseous phase where
they may form v h (l) mols (because of dis or association).
By approximation we can neglect the change (in the process) of the
volume of the condensed phase by comparison with that of the vapor.
The quantity AF in eq. (9.15) denotes, then, simply the increase of
volume of the gaseous phase due to VK (I) additional mols of the com
ponent h. According to eq. (1.18), this volume, before the beginning
of the process, is V = (W (1) + . . . + N a ( ")RT/p\ after it takes place,
the mol number Nh (l) is increased by VH (I} , the difference is, therefore,
AF = VK^RT/p. (9.25)
Of course, we suppose that N h (l) is very large compared with v h (l)
so that there is no appreciable change in concentration during the
process. (Or else the mol number changes are not the integers Vh
themselves but infinitesimal numbers 5N h , proportional to v hl as in
section 42). From the substitution into eq. (9.15)
v <>
p
a relation which can be integrated. Since T is regarded as a constant
in the partial derivative, the constant of integration with respect to p
will be independent of p but, in general, not of the temperature T:
log K =  Vh log p + v h log k k (T), (9.27)
where the second term represents the constant (or, rather, function
of T).
Applying this to the special case of eq. (9.20), with v h (l) = 1, we find
If we denote by p h (1) the partial pressure of the component h in the
gaseous phase, we can write # A (1) = p h (l) /p, according to eq. (8.06),
whence
x k = ph (1) /k h . (9.29)
When the mol fraction of a component in a liquid (dilute or perfect)
solution is varied at constant temperature, the partial pressure of its
vapor varies in the same proportion, and vice versa. In particular, we
can obtain from this equation some information about the solubility
1X62 DILUTE SOLUTIONS 161
of permanent gases in water and in other liquids: with neglect of quan
tities of the second order in N^/No, the amount of the gas absorbed by
the solvent is proportional to #* (l) and, therefore, proportional to the
partial pressure of the gas in the atmosphere over the solution. This
law was established empirically by the English chemist William Henry
(1803) and, accordingly, is known as Henry's law.
The amount of the absorbed gas (denoted anph) is usually expressed
in terms of the volume which it would occupy at C and 760 mm Hg.
The quantity a h is then called Bunsen's coefficient of absorption (if p is
given in mm Hg). The constancy of the coefficient a can be judged
from the following data referring to the absorption of nitrogen (N2)
in water at 25 C. 1
TABLE 21
HENRY'S LAW
/>Ar 2 (mm) 270 300 400 500 601.6 700 800 830
14.38 14.37 14.31 14.26 14.20 14.28 14.36 14.38
The trend of the dependence of kh upon T can be deduced from
eq. (9.16). Substituting the expression (9.27) into it, we find
d log k h Q
~
The latent heat Q of driving out an absorbed gas from the solution
is positive. Therefore, kh increases as the temperature is raised, so
that solvents absorb gases and vapors better when they are cold than
when they are heated. As an example we give the values of k for
(argonfree) nitrogen dissolved in water. 2
TABLE 22
tC 5 10 15 20 25 30 35 40 45 50
X10 7 4.084.57 5.07 5.55 6.00 6.43 6.85 7.23 7.61 7.99 8.37
In the general case, when the possibility of association (or dissocia
tion) is taken into account, eqs. (9.19) and (9.27) give
*A = (kk/P)*; p h ( "/x*' OK = ** (9.31)
A case in point is the solubility of hydrogen in molten copper
(at / = 1123 C). Let M be the mass of hydrogen, in mg, dissolved
in 100 g of copper. 3
1 Drucker and Moles, Zs. phys. Chemie 75, p. 405, 1911.
2 1. C. T. from measurements by Fox (Trans. Faraday Soc. 5, p. 68, 1909).
A. Sieverts and W. Krumbhaar, Zs. phys. Chemie 74, p. 294, 1910.
162 TEXTBOOK OF THERMODYNAMICS 1X63
TABLE 23
/>(mm).... 281 403 606 775 883 971 1046
M ......... 0.380 0.443 0.549 0.610 0.680 0.705 0.745
p*/M .... 43.5 44.2 43.,7 45.1 45.1 45.3 44.2
Hence g =  or M = iM (1) the hydrogen dissolved in copper must
be atomic. The constancy of p*/M has been ascertained for other
gases dissolved in molten metals (62 in Ag, H2 in Ni and Fe, etc.). It is
in keeping with this that the absorption increases with temperature.
According to (9.30) this means that the process of driving the gas out
of the metal is exothermic.
In many organic solvents we find association of the gas molecules.
However, it is usually incomplete, leading to irrational values of g*.
Exercise 88. From the data about nitrogen in water and from eq. (9.30) calcu
late the approximate heat of solution Q. (Per 1 mol, v^ = 1).
Exercise 89. Suppose that the same gaseous phase (1) is adjacent to two con
densed solvents. Equation (9.31) is then valid for either of the solvents. Show
that the two relations so obtained are compatible with Nernst's law of distribution
for the two condensed phases.
63. Raoult's law. The symbol p in eq. (9.31) represents the joint
pressure of the liquid solution and of its vapor. According to our dis
cussion in section 41, it is completely determined when the composition
of the system and its temperature are given. It is called, therefore, the
boiling pressure of the solution under these conditions. Equation
(9.31) is true for every component: let us apply it to the solvent
(h = 0) of a dilute solution. (As in the preceding section the con
densed phase is " preferred/ 1 we shall emphasize it by dropping its
superscript 2 while reserving the superscript 1 for the vapor phase).
rn  . <*>  MO/*> (I) ). (9.32)
This relation holds for every composition and, in particular, for the
limiting case of the pure solvent characterized by N\ =* . . . = N 9 =
and#o = #o (1) 1. The pressure corresponding to this limiting case is,
then, the boiling pressure p$ of the pure solvent. We can write, there
fore, 1 = PQ/HQ or po = feo whence
This relation takes a very simple form when the solutes are non
volatile (e.g. most salts) so that their partial pressures in the vapor
1X63 DILUTE SOLUTIONS 163
are negligible. The gaseous phase can, then, be regarded as pure
(*o (1) = 1)
p/po = so 17 ' . (9.34)
On the other hand, the mol fraction *o = No/N can be represented
as 1 (Ni+ . . . + N ff )/N, or by approximation
so _ 1 _(#! + ...+ NJ/No. (9.35)
Whence, within the accuracy of the present theory which neglects
terms of the second order in N h /No:
PO gO No
The presence of nonvolatile solutes lowers the vapor pressure of a
liquid. This fact was known to the chemists of the eighteenth cen
tury. In 1847 it was noticed by Babo that the relative lowering,
(p Po)/Po> is independent of the temperature. A few years later
(1858) Wiillner found it to be roughly proportional to the concentra
tion of the solution, and his observations were amplified by Ostwald.
However, the most extensive investigation of this subject is due to
Raoult (18861887), who established experimentally the law (9.36)
according to which the relative lowering of pressure is determined
solely by the mol fractions of the dissolved salts, being quite indepen
dent of the nature of the solvent and of the solutes. In earlier work
(1878) Raoult also formulated the laws according to which the boiling
and melting temperatures are changed by the presence of solutes.
They form the subject of the next section.
In the case of a single solute (1) eq. (9.36) can be written in either
of the two alternative forms
(9.37)
, m
go mipo 10
where m\ represents the number of mols of the solute in 1000 g of the
solvent, called its molality. The second form follows from the first by
substituting x\ = N\/No and noticing that in 1000 g of the solvent
noNo = 1000 and N\ = mi . In the case of aqueous solutions MO (I)S=: 18:
the quantity k m should, therefore, be numerically equal to 1.80. How
ever, in the tabulated data N\ is, of necessity, not the actual mol
number of the solute (which is usually unknown) but "the gram
formula weight", i.e. the mol number as it would be if the solute were
neither dissociated, associated, nor hydrated. The measured values
164
TEXTBOOK OF THERMODYNAMICS
1X63
of k m at low concentrations are for many substances close to 1.8, as
appears from Table 24.
TABLE 24
LOWERING OF THE BOILING PRESSURE
Solute
mi
km
Solute
mi
km
HsPO 4 ..
1
1 80
Citric acid
1
1 90
Mannitol
1
1 775
Pb(C2H 8 O 2 )2
1
1 87
Sucrose
10 3
1 82
H 3 AsO 4
1
1 93
Glycerol
0.5
1 78
In general, electrolytic solutes have tabulated coefficients k m larger
than 1.8 because they are often dissociated. Integral and halfintegral
multiples of 1.8 (K 2 S 2 O 3 3.65, Na 2 WO 4 3.64, Ba(NO 3 ) 2 3.55, etc,) are
fairly common in them. On the other hand, organic solutes show
mostly values of k m under 1.8, and this may be due to association.
For nonaqueous solutions, the data are usually given in the first
form (9.37). When go = 1 (the solvent does not associate), the
formula requires k x = 1. This requirement is satisfied in a large
number of cases, provided the concentration is sufficiently low, and
even for values of #, from 0.05 to 0.10, it is often still approximately
fulfilled.
TABLE 25
LOWERING OF THE BOILING PRESSURE
Solvent
Solute
Xi
k x
Benzene
Naphthalene
0.005
1.02
Ethyl alcohol
Benzil
0.001
0.99
Mercury *
Ag
0.0025
1.01
4 *
Tl
0.05
1.05
4 1
Zn
0.10
0.93
< 4
Au
0.05
0.55
*Hildebrand and Eastman, J. Am. Chem. Soc. 37, p. 2452, 1915.
Of course, here too, examples of association (and dissociation) of
the solute are not missing, as is indicated by the figure 0.55 relating to
gold : this metal must have in mercury the molecule Au2.
Exercise 90. In the last column of the following table are listed the measured
values of the relative lowering of vapor pressure (po  p)/p* per 1 g of solute in
1000 1 of solvent.
1X64 DILUTE SOLUTIONS 165
SOLVENT
SOLUTE
M. W
Water
Mannitol
18
i <
Glycerol
18
Br 2
BrI
179.8
Benzene
Naphthalene
78.05
t W(pQ  p)/pQ
20 C 0.978
1.935
8.70
4.92
Ethyl alcohol Benzil 46 . 05 15 2.14
Calculate the approximate molecular weights of the solutes from formula (9.37).
The accuracy is not very high.
64. Influence of solutes on the temperature and pressure oi
transformations. Let the temperature and pressure of equilibrium
of two phases of the pure solvent be TO, po. If small amounts of solutes
are added to the two phases, the equilibrium temperature and pressure
will be slightly changed, assuming the new values T, p but little dif
ferent from To, po The function log K depends only on p and T, and
we can expand it into the double Taylor series
log K(p, T) = log K(po, To) + (p po)
neglecting terms of higher order.
Equation (9.17) applied to the solvent (h = 0) gives
i>o (1) log *o (1)  "o (2) log * (2) = log K(p, T).
When the solvent is pure in both phases, the mol fr 'ty
#o (1) = # (2) = 1 and the left side vanishes, so that log f *'
For the partials of log K we can substitute from eqs. (9, ' 0^55
AF 'If/ weights of
Ri*$
1 7V/fc of eq. (9.43)
We shall restrict ourselves to the foil from these data the
vaporization of a liquid solution (2) conti  acid (MO 60.03).
utes. Second, freezing of a liquid solutic
not soluble in the solid solvent. In bot 1 b V the physiologist
in our formulas, will be pure (*o (1 > = T on a semipermeable
or freeze out while the solutes remain i , '' through the solvent
phase is the preferred one, in the sense " and is denoted by P.
we shall bring it into prominence b / dr ' er P a * of a cylindrical
"*. For simplicity, we
166 TEXTBOOK OF THERMODYNAMICS 1X64
0(2) SBJQ we wr ite, as in the preceding section, l (N\+ . . . +N v )/No t
whence
log *o   (Ni + . . . + NJ/No, log *o (1>  0. (9.40)
As we know from the theory of the phase rule (section 41), only
one of the two variables T and p can be chosen arbitrarily when the
composition of the liquid phase is given. We have, therefore, to dis
t nguish two cases in analyzing eq. (9.39).
(1) We set arbitrarily T = TQ and find the difference of the vapor
pressures of the solution and of the pure solvent, at the same tempera
ture To,
RTo Nl + ' + N *
The sign of the difference p po is opposite to that of the change
of volume AF in the transformation. In the case of vaporization, we
can use for A V the expression (9.25) and obtain with (9,18) the equation
(9 . 42)
(go = j>o (1> Ao) which is identical with (9.36).
(2) We set p = po and obtain the difference of the boiling points
f the solution and of the pure solvent, at the same pressure po. The
(accc Miing of the heat of reaction Q p is here the latent heat of vaporizing
Th, freezing VQ mols of the pure solvent; we shall denote it Q p = volo,
aratinr % is the latent heat referred to 1 mol of the liquid phase.
in the,
rtdy* W1T.+ +*
the solvef^ *o NO
where A V
1 mo i O f t j^ of vaporization the latent heat is positive, in that of
p&VdNo t>\* ^here exists, therefore, an elevation of the boiling
the solution wii '*' iing &* ni ' both due to the solutes  As
its contact with ' \ s much larger in absolute va lue, the
since its boiling pr, *o.^:derably less than the melting point.
its original position ^ e quantities N * in formula ( 9  43 ) re P r e
Po is  RTgndN 1 ' lumbers (taking into account association,
that dN mols of the r nd nOt the ^ ram " formula weights. The
then on, condenaatV computed for every solvent. Usually,
remainder of the ' ^ ^^^ ^ n ^^ ^ ^ s l ven t (normal
corresponds to 1 md^f numerical values for several solvents
the change of volur
1X65
DILUTE SOLUTIONS
TABLE 26
167
T B
B.P.
elevation
T M
M.P.
lowering
Water
3731
0521 C
2731
186C
Ethyl ether
307 6
2 18
1 79
Ethyl alcohol
351 6
1.17
Benzene
353 35
2.62
278 5
 5 10
Chloroform
334 3
3 82
209 6
4 80
Acetone
329.2
1.76
Acetic acid
391 2
3 11
289 7
3 90
Aniline . . ....
457 5
3 61
266 1
5 87
Phenol
455
3 54
314
7 4
Nitrobenzene
484.0
5 73
278 8
7
This effect offers a convenient and powerful method for the deter
mination of unknown molecular weights of solutes which was of great
help in the development of our chemical knowledge. The theoretical
treatment of these phenomena was due to Van t'Hoff.
Exercise 91. In the last column of the following table are given the measured
boilingpoint elevations corresponding to a concentration c of the solute in g pf
1000 g of solvent given in the third column
SOLVENT
Ethyl alcohol
Benzene
Acetone
Aniline
Phenol
Acetic acid
SOLUTE
I:
!
HgCla
Benzil
Benzil
Picric acid
c(g/1000g)
25.4
12.7
54.3
105.0
21.0
11.45
TTs
0?121
0.13^
0..*
i
Calculate from Table 26 the opproximate mr*
the solutes.
Exercise 92. What is the relation between t'
and the data of Table 26 which refer to molaliti
latent heats for water (MO =18), benzene (MO =
65. Osmotic pressure. It was dis
Pfeffer (1877) that solutes exercise 2
membrane which is impenetrable to t!
freely. This pressure is called osmoti
Suppose that the liquid solution fills
vessel (Fig. 21) closed above by the
168 TEXTBOOK OF THERMODYNAMICS 1X65
shall assume (at first) that the solutes are nonvolatile so that the
vapor, in the upper part of the vessel, is free from them; later we shall
drop this restriction. Suppose that we have in the vessel a second
piston consisting of a semipermeable membrane of the sort just men
tioned whose initial position coincides with the plane border surface
MN between the liquid and the gaseous phases. We push this piston
a little down, in a reversible isothermal way, to
the new position M'N', the volume between the
planes M'N' and MN being dV. As a result,
there will remain only pure solvent above the
membrane M'N': It is a process for recovering
a certain amount of solvent, say dNo mols, from
the solution. Since the volume V of the solution
is a function of TVo, N^ . . . N ff , while only NQ
changes, we have
FlG. 21. Calculation b
of osmotic pressure. dV = (dV/dNo)dNo = VodN ,
where #o is the partial molal volume of the solvent (compare section 32).
In pushing the membrane down, work must be done against the forces
of osmotic pressure P to the extent
DW = PdV= PVodNo, (9.44)
>rding to section 7).
re is, however, another reversible and isothermal way of sep
^'^Vo mols of the solvent from the solution. This can be done
* wing two steps: (1) By moving the upper piston up, infi
and at constant temperature, we vaporize dNo mols of
^he change of volume of the system is, then, AVdNo,
s the change corresponding to the vaporization of
. In this expansion the system does the work
r pnssure of the solution. (2) We cover
ly impermeable membrane. Having lost
the vapor is superheated at the pressure p
'> p). As we push the piston down into
done till the pressure assumes the value
iccording to eq. (3.32), and considering
^spond to godNv mols of the gas. From
the pressure stays constant, and the
ck is po&VodNo, where AFo again
vent. If it is permissible to neglect
id, compared with that of the vapor,
IX 65 DILUTE SOLUTIONS 169
and to use for the gaseous phase the laws of perfect gases, AF=oAFo:
the total work done in both steps of the process becomes
DW =  RTg Q dNo log (po/p). (9 AS)
In the beginning and the end of this process the system is precisely
in the same state as in the other method of separation carried out with
the semipermeable membrane. Moreover, both processes are isother
mal, therefore, DW' = DW, since both elements of work are equal to
the change of the function ^ between the initial and the final states
(compare section 34). Equating the expressions (9.44) and (9.45), we
obtain
g*. (9.46)
We have assumed, for simplicity, that the solutes are nonvolatile.
However, this restriction is not essential and can easily be dropped:
we need only imagine from the start a semipermeable membrane
(opaque to the solutes) at the surface of the solution. This device
permits us to have the pure vapor of the solvent in equilibrium with
a solution containing volatile solutes, since the vapors of the solutes
cannot penetrate the membrane. We can apply the reasoning leading
to eq. (9.46) to the general case. This equation remains valid with
the slight change that for p must be substituted the partial pressure
of the solvent in the gaseous phase pxv (l) :
(9.47)
The osmotic pressure depends on the ratio of the partial vapor pres
sures of the solvent, in its pure state and in the solution, being propor
tional to the logarithm of that ratio.
Equation (9.44) permits another interpretation of the osmotic
pressure: It has the form (2.02) of the element of work in the case of
generalized forces. The differential dNo represents the change in mol
number of the solution due to the removal of solvent. Pvo is, there
fore, the force with which the solution resists (because of the negative
sign) such a removal. In other words, it can also be regarded as a
tension or suction pulling the solvent back into the solution across the
semipermeable membrane. Pfeffer's method of measuring the osmotic
pressure is based on this interpretation (Fig. 22) : A glass tube, \xJiO8e
lower end is closed by a semipermeable membrane MN (made of cop
170
TEXTBOOK OF THERMODYNAMICS
1X65
M
B
H 2
FIG. 22. Measure
ment of osmotic
pressure.
per ferrocyanide), is partially immersed in a beaker filled with pure
water. When cane sugar is added to the water inside the tube, the
solution in it rises because it sucks in water from the beaker. The rise
stops when the pressure of the column AB is exactly equal to the force
of suction, i.e. to the osmotic pressure P. This pressure is by no
means small: In the case of more concentrated
solutions, it is better to connect the tube with a
manometer because the column becomes incon
veniently high. In fact, pressures up to 268 atm
have been measured in sugar solutions. The fol
lowing familiar demonstration experiment shows
how powerful the osmotic pressure can be. A part
of an ox bladder is formed into a bag, filled with
a strong aqueous solution of ethyl alcohol, and
tightly sealed. The bladder is semipermeable in
that it holds back alcohol but lets through water.
When the bag is immersed in a waterfilled vessel,
the water is drawn into it with such force that the
bladder cannot stand the strain and bursts within
a short time.
Let us turn to the general case when there are solutes on both sides
of the membrane permeable only for the solvent. Both solutions
exercise, then, a sucking action upon the solvent, and the resultant is
a tension equal to the difference of the osmotic pressures, directed
towards the side of higher concentration. Two solutions of the same
osmotic pressure leave the solvent in equilibrium and are called isos
motic or isotonic. On these facts is based a method for determining
the osmotic pressure within living cells, whose integuments are natural
semipermeable membranes. When a living cell is placed in an aqueous
salt solution of higher osmotic pressure than its own, it begins to lose
water and to shrink. When the osmotic pressure of the medium is
lower, it swells. Only in solutions approximately isosmotic with its
interior does it retain, under the microscope, its size and general appear
ance. By this method of finding a solution isosmotic with them, the
osmotic pressures in many physiological cells have been determined.
In plant cells they are sometimes as high as 14 atm. In general, the
vegetable and animal cells are nearly isosmotic with the natural saps
and animal fluids in which they live. For instance the red corpuscles
of the human blood and most other cells of the human body have the
same osmotic pressure as the normal blood serum, which in turn is
isosmotic with a solution consisting of 0.156 mol of NaCl per 1000 g of
water. The harmful effect of certain diseases is believed to be due to
1X66 DILUTE SOLUTIONS 171
the fact that they upset the osmotic equilibrium of the body fluids and
cause a destruction of cells similar to the breaking of the ox bladder
in the abovementioned experiment.
66. Van t'Hoff's equation for osmotic pressure. The expression
for osmotic pressure (9.47) takes a simple form in the case of dilute
and perfect solutions. On one hand, the partial molal volume then
becomes identical with the ordinary molal volume of the solvent:
t/o = VQ. On the other, we may substitute for o/p#o (1) from (9.33)
and obtain
p=^log*o, (9.48)
VQ
an expression which is valid both for volatile and nonvolatile solutes.
If the solution is dilute, log XQ takes the form (9.40) and the volume
NQ VQ of .the solvent can be replaced by the volume V of the solution
(V = NQVQ), whence
PV (Ni + . . . + N ff )RT. (9.49)
This is Van t'Hoff's equation for osmotic pressure, which is formally
identical with eq. (1.18) for a mixture of perfect gases. By specializa
tion, we can obtain from it the partial osmotic pressures of the indi
vidual solutes: if only the component h is present in the solution,
P h V  NiJRT, or P h x h RT/v Q . (9.50)
It follows from (9.49) that the partial osmotic pressures are addi
tive P = PI + . . . + P . In other words, the solutes in a dilute
solution behave like perfect gases.
The experimental data are usually expressed in terms of the
molality f i.e. mol number of the solute per 1000 g of the solvent, which
is denoted by m h . For aqueous solutions, this reduces to NQ VQ
1000 cm 3 and, together with eq. (9.50), gives P h  RTm h /lQQQ, or
numerically for C,
P k  22.4m*atm. (9.51)
A solution of the molality m = 1 is called normal. We see that
the normal solution has a theoretical osmotic pressure of 22.4 atm.
If the molality is computed from the gramformula weight, eq. (9.51)
can give a correct result only barring association, dissociation, and
hydration. A few of the experimental values for dilute aqueous solu
tions (directly measured with copper ferrocyanide semipermeable mem
branes) are given in Table 27. The values of mi listed in this table
are not corrected for association, etc.
172
TEXTBOOK OF THERMODYNAMICS
1X66
TABLE 27
OSMOTIC PRESSURES
Solute
mi
P (atm)
P/mi
Arnygdaline
0.0219
0.474
21.7
Antipyrine
0.0530
1.18
22.3
Resorcinol
0097
199
20 5
Saccharin
0.0029
070
24.3
Glucose
0.555
13.2
23.8
Mannitol . ...
549
13 1
23.9
When the solution is not dilute we have to fall back on the general
eq. (9.48). For aqueous solutions, vo = 18.015 and the coefficient
has the numerical value (at C)
P =  2865 logio * atm. (9.52)
This formula can be tested for sucrose. 1 It is known that sucrose
forms a hydrate, each molecule attaching five molecules of water. The
molality of pure water is 1000/18.015 = 55.51; consequently, in the
sucrose solution it is wo = 55.51 5wi and l/#o = l+wi/(55.51 5wi).
The values thus obtained are listed in Table 28.
TABLE 28
OSMOTIC PRESSURE OF SUCROSE
P (calc.) . . .
P (obs.) . . .
0.1
2.24
2.25
0.3
6.90
6.91
0.6
14.12
14.22
1.0
24.4
24.76
2.0
53.5
54.9
3.0
89.0
90.0
4.0
132.5
129.5
The agreement is surprisingly good and shows that the sucrose solu
tion can be regarded as perfect, as far as this calculation is concerned.
\ In a perfect solution the difference between solvent and solute dis
appears. It is better, therefore, to avoid the word " solvent " and
to speak instead of the " component " without making any restric
tions as to its mol number. For a binary solution the interpretation
of the preceding section can, then, be stated as follows: the partial
osmotic pressure PI represents the force with which the other component
is held in the solution (and vice versa). According to (9.48) it is
proportional to the logarithm of the mol number of that component.
If we gradually remove it from the solution, by the mechanism of
1 Lord Berkeley and Hartley, Proc. Roy. Soc. (A) 92, p. 477, 1916; H. L. Cal
lendar, Proc. Roy. Soc. (A) 80, p. 466, 1908.
1X66 DILUTE SOLUTIONS 173
Fig. 21 or by some other method, the difficulty of the removal con
tinually increases and becomes prohibitive in the end, because the last
traces are held with an infinite force.
This impossibility of removing the last vestiges of any component
from a phase is quite general, as can be seen, in the simplest way,
directly from the fundamental equation (9.17) which applies to this
case as the solution is then certainly dilute with respect to this com
ponent. Log K is, by its definition (9.14), a finite number: therefore,
Xh (l) cannot become zero. Otherwise the left side of eq. (9.17) would
become infinite, unless J0 (2) were zero at the same time. This means
that a component cannot be completely absent from one phase of a
system if it has a finite concentration in another phase. All parts of
our earth have been in direct or indirect interaction for immense
periods of time. Consequently, every sample of terrestrial material
must contain minute quantities of all known substances, and an abso
lutely complete purification of it is impossible. However, the second
interpretation of the osmotic pressure here used should not obscure
the fact that it is not an abstraction, but, in every way, a real pressure
like that of a gas. After all, the only tangible thing about any pres
sure is that it is able to resist the motion of a piston or to set it in
motion, and there is no difference between the osmotic and the ordi
nary pressure in this respect. We shall gain some insight into the
deeper nature of the osmotic pressure in section 146 but shall bring out
in this chapter the analogies between dilute solutions and perfect gases.
We have already seen that the osmotic pressure has a tendency towards
uniformity: if the concentration of a solution is nonuniform, the local
differences of osmotic pressure set up diffusion currents which make
it uniform (barring gravitational fields, compare section 108). In
this process some mass of the solute (K) may be transferred from a
place with the osmotic pressure Ph to one with P' h (in an isothermal
way). The work done in this transfer must have in a dilute solution
the same expression (3.32) as in a perfect gas, in view of the identity
of the equations of state. Thus we obtain (per mol of the solute)
w = RT log (Pk/P'd = RT log (x k /x' k ). (9.53)
Exercise 93. What are the osmotic pressures of the following solutions, sup
posing that they satisfy the formula (9.51): (a) antipyrine (CiiHi 2 N 2 O, n 188.11)
10 g in 1000 g of water; (b) saccharin (CyHfiNOaS, M = 183.1 1), 2 g in 1000 g of water;
(c) glucose (CHi 2 O, /* = 180.09) 100 g in 1000 g of water?
Exercise 94. What are the osmotic pressures of sucrose solutions (CuHtsOu,
M = 342.17), according to the formula (9.52) and taking into account hydration aa
in Table 28: (a) 250 g, (b) 500 g, (c) 1000 g in 1000 g of water?
174
TEXTBOOK OF THERMODYNAMICS
1X67
Exercise 95. Calculate numerically from eq. (9.53) the work (in calories) that
must be expended to reduce the mol fraction of a solvent in the ratio 1:10, at 20 C.
67. Saturated solutions. Solubility. In general, a solvent can
take up only a limited amount of a solute at a given temperature. For
instance, when we gradually add a salt or other solid solute to water,
it dissolves at first, increasing the strength of the solution. However,
when a certain limiting concentration is reached, further additions do
not go into the solution but form a solid precipitate at the bottom of
the vessel. Such a solution is called saturated, and the solute in it is
in equilibrium with its solid crystalline phase. The behavior of liquid
solutes is analogous, but we shall restrict ourselves in this section to
solids, postponing the discussion of liquids until the next chapter.
The mol fraction x, of the solute in the saturated state is sometimes
called its solubility. More often, however, the solubility is defined as
the molality m, of the solute in the saturated solution. If we denote by
mo the molality of the solvent (mo = 1000/Mo),
, = w,/(mo + m,).
(9.54)
(For water mo = 55.51). The solubilities of some of the more common
solutes in water, at C, are as follows:
TABLE 29
SOLUBILITIES
Solute
m.
Solute
m,
Solute
m.
NaCl
6.10
KClOs
270
CuSO 4 .
88
NaHCOs
0.82
CaCh
5.35
FeSO 4
1 030
NaOH ...
10 50
Ca(OH) 3 .
0239
Sucrose
5 36
KC1
3.76
CaSO 4
01292
Lactose . . .
345
KOH
17.3 *
AgNO 8
6 65
Maltose . .
1 70
It should be mentioned that the table is meant to illustrate merely
the general notion of saturation. The substances listed in it are for
the most part " strong electrolytes ", they do not obey quantitatively
the theory of dilute solutions of this section but need a correction which
will be given in section 115, Moreover, the numbers m, listed in it
present the " gramformula weights " of the undissociated and unhy
drated solutes.
1X67 DILUTE SOLUTIONS 175
Returning to the theory, we shall assume that the solvent is not
soluble in the solute (as is the case with most solids) so that the excess
of the solute represents a pure phase (xh (l) = 1). In applying to this
case of equilibrium the fundamental eq. (9.17), we label the excess
solute (1) and drop the superscript for the liquid solution
^log*M = logK. (9.5S)
Both phases are here condensed, so that the change of volume in
the process of solution is small and may be neglected in the first
approximation (AF = 0). According to eq. (9.15), this means that K
is independent of the pressure and is a function of the temperature
alone: K = K(T). At the same time the heat, Q p = Qv, developed in
the process is entirely due to the difference of the internal energies in
the two states Q p = A 7 because, within this approximation, the work
done by the system is negligible. If we denote by Aw* the difference
of internal energies per 1 mol of solute, Q p = ^Aw*, and we find, from
eq. (9.16)
d log X* Aw*
(9 ' 56)
or substituting for Xk the expression (9.50), we obtain
d log P h _ Aw* + RT _ l h
dT ~ RT* ~~RT*
(9.57)
This relation is formally identical with the approximation (7.07) of
the ClapeyronClausius equation. The process of solution of a salt is
completely analogous to the process of vaporization, and the osmotic
pressure PK of the saturated solute corresponds to the boiling pressure
ps of the vapor. In fact, the quantity k in eq. (9.57) has the same
structure as the heat of vaporization: According to (3.21) and (3.24),
/ = Aw + p&v, or, when the volume of the condensed phase is neg
lected and the vapor regarded as a perfect gas, / = Aw + RT.
The heat of solution Aw* is almost invariably positive. Therefore,
we conclude from eq. (9.56) that the solubility of a salt increases with
temperature, unlike the solubility of a gas (compare section 62). Our
assumption that the volume does not change is only a first approxima
tion, in reality AF has, usually, a very small but finite value. Conse
quently the effect of added pressure consists in a very slight increase or
decrease of solubility.
176 TEXTBOOK OF THERMODYNAMICS 1X68
Exercise 96. The solubility of many aqueous solutes is well represented by the
formula
1 0.052234
logio +B.
FORMIC ACID ACETIC ACID GLYCEROL MANNITOL
Range.. 7.Oto8.4C 27 to 10 1.3 to 18 40 to 100
A 12090 14980 17150 2142
B 2.243 2.70 3.078 2.102
Calculate from (9.56) the approximate heat of solution AA within the range, assuming
that there is no dissociation or hydration.
68. Chemical equilibrium of solutes. We suppose that chemical
reactions are possible between the components of a dilute solution,
according to eq. (6.46)
*>oGo + vid + . . . + v 9 G 9 = 0. (9.58)
The general equation of equilibrium (9.12) takes then the form
VQ log #o + v\ log x\ + . . . + v 9 log x ff = log K. (9.59)
Since we restrict ourselves to dilute solutions, the volume is prac
tically determined by the amount of the solvent, and its change in the
reaction (9.58) is very small. As in the preceding section, we conclude
from this that K is very little dependent on the volume and with a
good approximation can be regarded as a function of the temperature
alone, K = K(T). Moreover, we can replace x h = N h /N by N h /No,
because the difference would amount only to terms of the second order
in the expression (9.59) which are systematically neglected in the
theory of dilute solutions. In application to the solvent, this means
that we can replace XQ by 1, obtaining from the last equation
siW... x." = K(T), (9.60)
without regard to whether the solvent takes part in the reaction or
not. Because of the relations (9.50) this is equivalent to
P^Pa* 1 . . . P/ r = K P (T) = K[), (9.61)
\ VQ /
(v = v\ + V2 + . . + O
This equation is formally identical with eq. (8.17) for a mixture of
perfect gases. However, the conditions under which the equilibrium
takes place are different in the two cases: The mixture of perfect gases
is usually observed at a given pressure p which is an important factor
1X68 DILUTE SOLUTIONS 177
in determining the partial pressures ph, when only the mol numbers
are known. On the other hand, the partial osmotic pressures P* of the
solutes are completely determined by N h /No = WA/WO and have noth
ing to do with the pressure. This is brought out clearly, if we write
the condition in the third form, resulting from (9.60)
mi'W 1 . . . w/' K m = K(r).m v (9.62)
(We recall that WQ is the molality of the solvent: wo = 1000 /Vo)
The relation between the molalities depends only on the temperature
and on the atomic weight of the solvent.
As in the case of gases, the most interesting question is the deter
mination of the degree of dissociation. Let us take as an example the
same reaction (8.22) which we discussed in section 52, the dissociation
of nitrogen tetroxide (dissolved in chloroform). The expressions for
the numbers VH and Nh remain the same as those given there. Sub
stituting them into (9.60), we obtain for the degrees of dissociation
2 K**. (9.63)
where m\ is the molality of N2CU originally dissolved.
Measurements on the dissociation of N2O4 in chloroform were car
ried out by Cundall. 1 Lewis and Randall 2 give the following table
calculated from his data (for C).
TABLE 30
CHEMICAL EQUILIBIRUM OF NITROGEN TETROXIDE
#N204 #N02 10 &
0.018 0.00010 49
0.037 0.00012 35
0.050 0.00015 43
0.066 0.00019 52
0.125 0.00029 66
The constancy of K is not very good owing, perhaps, to the insuffi
cient accuracy of the colorimetric determinations of # N02 . Of course,
the degree of dissociation is quite different from that which gaseous
nitrogen tetroxide has under the same conditions: KP comes out here
about 2.00 X 10" 4 atm, while (8.25) gives 0.020 atm for the gas
reaction at C. This is not surprising since the heats of reaction are
1 Cundall, J. Chem. Soc. 50, 1076, 1891; 67, p. 794, 1895.
1 Lewis and Randall, Thermodynamics, p. 303, New York, 1923.
178 TEXTBOOK OF THERMODYNAMICS 1X68
very different in the two cases: in the chloroform solution there must
be added the difference of the heats of solution of the two components.
For very high dilutions (mi/wo 1), approaches unity. This
is a general property of dilute solutions, as we can readily see from
eq. (9.64): Whenever the dissociation leads to an increase in the
number of molecules (v > 0), dilution (i. e. increasing NQ) has the same
effect as increasing K (compare section 52). It advances the disso
ciation which becomes complete in the limiting case of extreme dilution.
This influence of the relative amount of the solvent upon the degree
of dissociation is the analogue of the effect of neutral gases upon a gas
mixture discussed in section 54. The addition of neutral solutes has,
according to eqs. (9.61) and (9.62), no appreciable effect upon the
equilibrium. The reason for this is that in a dilute solution the quan
tity of a solute is, necessarily, small and its contribution to the avail
able volume negligible. On the other hand, the remarks made in the
sections 53 and 55 about the depressing influence of the excess of one
active component and about the mutual interdependence of two simul
taneous reactions remain fully valid in the case of solutions.
Exercise 97. Calculate the factors RT/v Q and w entering into KP and K m
for the following solvents: water (MO 18, density d = 1.00), chloroform
(MO  119.38, d  1.526), ethyl ether (/*o  74.08, d  0.736), benzene (MO = 78.05
d  0.900).
Exercise 98. Carbonic acid and acetic acid in aqueous solutions dissociate into
ions according to the following equations:
H+ + HCOa"  H 2 CO, (aq) =0, K m  1.84 X lO" 6 ,
H+ + C 2 H 8 Or  C 2 H 4 O 2 (aq)  0, ** 3.50 X 10' 7 .
Calculate the degrees of dissociation for the molalities mi (of the acid originally
dissolved) 0.001/0.01, 0.05, 0.10, 0.20.
Exercise 98'. Show that the equation of equilibrium (9.62) can be also written
in the form
' (9.64)
CHAPTER X
EQUILIBRIUM OF BINARY SYSTEMS 1
69. General remarks. The character of the equilibrium between
different phases of a system with two independent components
(binary systems) is of great importance in chemistry and metallurgy.
A vast body of experimental data, pertaining to it, has been accumu
lated in these branches of science. The role of thermodynamics is to
help in the explanation and classification of this material by enumerat
ing the theoretically possible types of equilibrium. Of necessity, much
of the information derived from thermodynamics is qualitative: The
problem consists in treating the equilibrium between solid, liquid, and
gaseous binary solutions (in the sense of section 56), but we do not*
possess the theoretical knowledge of the condensed phases which
would be necessary for a quantitative treatment, viz. the equations of
state and the energy expressions for the solid and liquid state.
Fortunately there exist, if not real, at least ideal systems in which the
theory can be completely carried through, namely, the simplified
models which we called perfect solutions in section 56.
The essence of this approximation is that the interactions of the
components are taken into account only in so far as they affect the
mutual solubility. Apart from this, each component is treated as if
it were alone. We propose, therefore, to work out, for every type of
equilibrium, the formulas and curves pertaining to the case of perfect
solutions. Comparing them with the experimental curves, we shall
see that this model is not oversimplified and gives a good account of
the main features of real systems, giving examples of every type of
diagram that occurs in experiments.
According to the discussion of section 41, the maximum number
of phases which can be simultaneously in equilibrium exceeds by two
the number of independent components: for binary systems it is 4.
Four phases can coexist only at the " quadruple point " for which all
the variables of the system (viz. temperature, pressure, and composi
tion of the phases) are perfectly determinate. In the case of the
1 This chapter may be skipped without loss of continuity.
179
180 TEXTBOOK OF THERMODYNAMICS X70
coexistence of three phases the system has one degree of freedom; i.e.
one of the variables can be chosen arbitrarily. For experimental rea
sons, this is, usually, the pressure : when p is given, the whole system
is determined, so that equilibrium can take place only at one definite
temperature and for definite compositions of all the three phases.
The most interesting case is, however, that of the coexistence of two
phases, because it was shown in the general theory of Chapter VI that
the other cases can be reduced to it. Then, two of the variables can
be chosen at random: for instance, we can prescribe the temperature
and the composition of one phase, so that equilibrium is possible at
only one definite pressure. Or else, we can select arbitrarily the pres
sure and the composition of one of the phases, and so completely
determine the equilibrium temperature.
The parameter which determines the composition of a phase is the
mol fraction of either component. If we indicate the components by
the subscripts 1 and 2, and the phase by the superscript (/), we find
from the definition (8.05) of the mol fraction
Xl <* + * a < = 1. (10.01)
In most of the applications of this chapter we shall suppose that
the components do not interact chemically. In this case, there is no
difference between the independent components and the actual
constituents.
A few remarks about ternary systems will be made in section 75.
70. Vapor pressure of binary systems at constant temperature.
The available experimental data have reference to the equilibrium of
liquid binary solutions and their vapors, because the vapor pressures
of solid solutions are too low to be conveniently measurable. There
fore, we shall now restrict ourselves to the liquidgaseous equilibrium
and we shall further assume that, in the liquid phase, the components
are mutually soluble in all proportions. We shall indicate the mol
numbers of the two components, in the liquid phase, by x\, #2 and, in
the vapor, by x'i, #'2, and shall apply the relations derived in the pre
ceding chapter for perfect solutions.
In a perfect solution either component can be regarded as the
solvent and obeys Raoult's law. If the molecules remain unchanged
in the process of vaporization, this law must be used in the form
(9.33), with go = 1,
px'i = PBI xi, px f 2 = pB2 x 2t (10.02)
where PBI and pB2 denote the boiling pressures of the two components
X70
EQUILIBRIUM OF BINARY SYSTEMS
181
in their pure state. Adding these two equations and taking into
account the relations (10.01) we find
P =
+
(10.03)
The plot of the pressure against the mol number x\ in the liquid
phase is a straight line (Fig. 23).
We can express the pressure also in terms of #'i, the mol number in
the gaseous phase. Dividing the eqs. (10.02) by psi, pB2 respectively,
and then adding them,
pBlpB2
P '
pBl(l 
(10.04)
This corresponds to the curve of hyperbolic character labeled
11 vapor " in Fig. 23. This figure represents the socalled pxdiagram:
i.o
1.0
1.0
FIG. 23.
FIG. 24.
FIG. 25.
Boiling pressure as a function of composition.
when the composition and pressure of the liquid phase are given by
the point A (on the " liquidus " curve), the constitution of the vapor
in equilibrium with it is obtained by drawing a horizontal to the point
B of intersection with the "vapor" curve.
When we turn to the experimental ^diagrams of pairs of organic
and inorganic liquids, such a large proportion of them is found to be
of the general type of Fig. 23 that it is justifiable to call it the normal
type. It is true, the representation is not quantitative in that the
" liquidus " locus geometricus is, generally, not straight but also
curved. But it is typical in that both curves are monotonic and the
vapor pressure of the solution intermediate to the boiling pressures of
the pure components. Especially, pairs of chemically closely related
substances (benzenetoluene, hexaneoctane, methyl and ethyl alcohol,
etc.) belong to this type. We have mentioned in section 56 that such
pairs must be expected to approximate the behavior of perfect solu
tions.
182 TEXTBOOK OF THERMODYNAMICS X70
However, there exist two other types of (p, jc)diagrams which are
illustrated by the Figs. 24 and 25. We shall call them the mrtype
and the mmtype. It is often somewhat uncertain what the molecular
weight in the liquid phase is. Therefore, the pressure is usually plotted
not against the mol fraction but against the mass fraction
yi  Mi/(Mi + M 2 ) = *i/(*i + /*2*2/Mi), (10.05)
Mi, M2 being the masses of the two components and /*i, /*2 their
molecular weights. This way of plotting has also some theoretical
advantages as we shall see in section 74.
Our simplified model (the perfect solutions) permits us to predict
a behavior of this kind in certain cases, viz. when the molecular
weights of the components are different in the liquid and gaseous states.
The generalization (9.33) of Raoult's equation must, then, be. used,
and this leads, instead of (10.03), to the equation
P = PBI xi l/ai + p B *(l  *i) 1/<7a , (10.06)
where gi = MI/M'I and #2 = ^2/^2 It is to be expected that the case,
when both components are subject to association or dissociation of
molecules, is a rare one. It will be sufficient, therefore, to assume that
only the exponent g\ is different from 1, while g% = 1. It is easy to
see that, under this assumption, the slope dp/dxi will, in general, go
to zero for a value of x\ between and 1, provided the boiling pressures
PBI and pB2 are of the same order of magnitude. This means that the
pressure reaches in this interval an extreme; whether it is a maximum
or a minimum depends upon the sign of the second derivative
d?p/dxi 2 (1 gi)gi~ 2 si#i (1 ~ 2ai)/w . In the case of association in
the liquid phase gi > 1, and the sign is negative; in the case of
dissociation, g\ < 1, and it is positive. We should, therefore, expect
the maxtype when the molecules of one of the liquid components are
associated, and the />mintype when they are dissociated.
On the whole, these expectations are borne out by the facts: It is
known that hydroxylated liquids have a tendency towards association
when mixed with another liquid which is free of the hydroxyl group.
Upon looking through the list of systems showing the m*xtype, we
find that it consists largely (although not exclusively) of such pairs of
liquids. On the other hand, aqueous solutions of HI, HBr, HC1, HF
which are considerably dissociated belong to the pmmtype. However,
there is another cause which may produce these kinds of pardiagrams:
the existence of chemical compounds of the two components. If the va
por pressure of the compound is higher than of either pure component,
X71 EQUILIBRIUM OF BINARY SYSTEMS 183
the pxcurve must have a maximum, if it is lower, a minimum. The
reason why the two curves in the Figs. 24 and 25 have a, joint maximum
or minimum will become clear in section 74.
This theory has an important practical application in the process of
distillation. In the normal pressure curve (Fig. 26) there corresponds
to the composition A of the liquid the composition B of the vapor;
that is, the vapor is richer in the component (2) with the higher boiling
pressure. Prolonged boiling at constant temperature, therefore, causes
the liquid to become poorer in this component so that its representative
point moves downward towards psi If the purpose of distillation is
to drive out the more volatile component, one has only to continue
boiling and, eventually, the liquid will consist of the practically pure
component (1). On the other hand, the pro
cedure is different when it is desired to prepare
the pure component (2). Starting again from the
state A, we let the solution boil for a time and
collect its vapor which has the constitution B.
Condensing this vapor we obtain a liquid of the
same constitution represented in our diagram
by the point C. Boiling this fraction (at the same
temperature) gives a vapor of the composition Z>, w >
etc. By such steps it is possible to approach FIG. 26. Distillation.
closely the point pB2 representing the pure volatile
component. It is clear that pressure curves of the mintype (Fig.
25) preclude the success of the first kind of distillation, and curves
of the maxtype (Fig. 24) that of the second. The bestknown
example is the waterethyl alcohol solution having a maximum at
4.43% water by weight so that the purification stops with 95.57%
alcohol (about 194 proof). Distillers have learned, in recent years,
to get around this maximum by adding a third component to the
solution.
71. Temperature of transformation at constant pressure. When
the pressure is kept constant, the temperature of equilibrium of two
binary phases is completely determined by the composition of one
of them. To treat this case theoretically for perfect solutions, we have
to fall back again upon the fundamental eqs. (9.17) or (9.22), but to
evaluate the equilibrium constant K in a way somewhat different from
that of the preceding section. We can apply to the first component
the eq. (9.16) in the form
, 10(m
(10 ' 07)
184 TEXTBOOK OF THERMODYNAMICS X71
where l\ is the latent heat of transformation referred to 1 mol of the
first component in the primed phase (v'i = 1, vi either integral or frac
tional).
We have pointed out in section 45 that the latent heat is a rather
slowly varying function of T which can be regarded as constant in the
first approximation. This approximation is amply sufficient for the
purpose of the present investigation because we have already intro
duced an element of inaccuracy by supposing the phases to be perfect
solutions. Integrating, we find therefore
log Xi =  h/RT + Ci(/>), (10.08)
and a similar relation with the subscripts 2, for the second
component. In the particular case, when the second component
is absent (xi = x'\ = 1), we find from (9.17) log K\ 0, and at the
same time the temperature of equilibrium is the transformation tem
perature Tbi of the pure first substance. This serves to determine the
constant as Ci(p) = h/RToi. There follows
h 1
(10.09)
On the other hand, the fundamental equations take the form (9.22),
with v\ = j/2 = 1,
Oc'i/^i) 17 " 1 = xi, (x'z/Kz) 11 ** = * 2 . (10.10)
We shall treat in this section the case when the components in both
phases ar^jnutually soluble in all proportions. Both components are
then present in either phase, and eqs. (10.10) must be treated as simul
taneous. The relations (10.01) permit us to eliminate all the variables
but xi. In particular, when there is no association in either component
(vi = V2 = 1),
*! =  (10.11)
This equation and x'i = K\x\ resulting from (10.10) give the
explicit expressions of the compositions of both phases (x\ and #'i) as
functions of temperature. They apply to all cases of twophase equi
librium, of which the more important (because of the accumulated
experimental material) are liquidvapor and solidliquid. The graph
ical representation of these formulas leads to the socalled (Tx)
X71
EQUILIBRIUM OF BINARY SYSTEMS
185
2032
exp.
theor.
diagram (or Jydiagram, if the temperature of transformation is
plotted against the mass fraction y). As an example we give in Fig. 27
the yTcurves computed with the help of our equations for the solid
liquid equilibrium of an alloy of platinum (Toi = 2032, /i = 5270
cal mol 1 ) and gold (7o 2 = 1338, / 2 = 3140 cal mol" 1 ) as dashed
lines. For comparison the experimental curves are drawn as solid
lines. The agreement is by no means quantitative, but the general
trend is well accounted for by the theory. It will be noticed that
the order of the curves is inverted by comparison with Fig. 23, the curve
giving the composition of the phase, stable at higher temperatures,
lying above the other. Apart from this, there is a complete analogy
between the cases of constant
temperature (preceding section)
and constant pressure inasmuch
as there exist, in addition to the
" normal" or monotonic type of
Fig. 27, also r max  and r m i n types
of curves (of the same shape as
those of Figs. 24 and 25). They
are, especially, a frequent occur
rence in the liquidvapor equi
librium of binary systems com
posed of organic substances. It
would be easy to generalize our
formulas for the case of any
numbers v and to show that association or dissociation in the liquid
phase leads to these types. However, this is unnecessary because
in section 74 we shall give the general proof that conditions which
produce the /wtype of the ^diagram also cause the r m i n type
of the TVgraph. In the same relation to each other stand the
mln and r max types. We can, therefore, refer, with respect to the
influence of association and dissociation, to the discussion at the end
of the preceding section.
Curves of r m i n type occur also in the equilibrium of binary metal
alloys with their melts, especially, when one of the components belongs
to the ferromagnetic group. We did not find in literature any discus
sion of the causes responsible for the minimum, but the prevailing
tacit assumption seems to be that it is due to the existence of inter
metallic compounds. It should be mentioned, however, that in the
case of CoMn alloys, a close investigation in Tammann's laboratory
revealed that compounds of these metals do not exist. 1 In view of
1 Hiege, Zs. anorgan. Chemie 83, p. 253, 1913.
0.0 0.5 1,0
FIG. 27. Temperature of fusion as a
function of composition.
186
TEXTBOOK OF THERMODYNAMICS
X72
this, we calculated from our formulas the Incurves (Fig. 29) which
would result on the assumption that Co (T<>2 = 1762, /2 = 3950
cal mol"" 1 ) has in the solid state a diatomic molecule when in alloy with
Mn (Tbi = 1523, k = 3560 cal mol 1 ). Comparing the theoretical
curves with the experimental 7>diagram (Fig. 28) by Hiege, 1 we see
that the minimum occurs at the right mass fraction (30% Co) but at a
temperature which is some 80 lower than that measured. Although
this discrepancy lies well within the limits of the theoretical inaccura
cies, we do not feel that the evidence is strong enough to prove posi
tively the existence of association, especially as it is not quite clear
what this would mean in terms of the lattice structure of the alloy.
1762^
1423
Co
1762,
1.0
0.5
FIG. 28.
Temperature of fusion as a function of composition
0.5
FIG. 29.
1.0
But we do wish to point out that there lies here a problem which, pre
sumably, could be settled by experimental (Xray) methods.
72. Case of mutually insoluble components. Let us suppose that
in one of the phases the components are mutually insoluble. This can
be illustrated by many examples of which we mention only a few : (1)
Aqueous solution of a salt which is insoluble in ice. When the liquid
solution is cooled, either a part of the salt is precipitated or a part of
the water congealed, conditional upon the concentration of the solu
tion. There are, therefore, two cases of two phase equilibriums:
liquid solution solid salt and liquid solution pure ice. (2) Molten
alloy of two metals which do not form mixed crystals. Upon cooling,
either of the metals can freeze out partially, according to the composi
tion of the liquid phase. (3) Mixture of the vapors of two immiscible
liquids. As the temperature is lowered, either one or the other liquid
condenses.
In all these cases we have equilibria of two kinds of pure substances
with the (liquid or gaseous) solution. If we suppose that the pressure
p is kept constant, eqs. (10.09) and (10.10) of the preceding section
1 Hiege, loc. cit.
X72
EQUILIBRIUM OF BINARY SYSTEMS
187
still apply, but we must put in them x\ = 1 or 002 = 1, the condition
that the unprimed phase is pure. Equations (10.10) become
x'l =
= (1 
= K 2 ,
(10.12)
and they are no longer simultaneous but refer separately to the two
possible kinds of equilibrium.
To make this quite clear, we shall consider a special example, the
molten alloy of cadmium (Toi = 596, /i = 1245 cal) and bismuth
(r 02 = 546, / 2 = 2110 cal). The first of eqs. (10.12) gives then the
composition of the liquid phase when it is in equilibrium with solid
bismuth; the second, when it is in contact with solid cadmium. Com
M
600
N
>>&#*
300
*Cd
0.55 1.0
FIG. 30. Solid lines experimental,
dashed and dotted curves theoretical.
* l.o
FIG. 31. Experimental liquidus
puting Ki and K% from eqs. (10.9), we find, for the first case, the theo
retical curve of A A' of Fig. 30; for the second, the curve BB'. Let us
start from a composition (x'\) and temperature (T) of the liquid alloy
represented by the point M. When this system is slowly cooled, the
representative point first moves down vertically, until it hits the
curve AA f . From then on, bismuth begins to freeze out; accordingly,
the cadmium concentration of the liquid phase increases and, as the
cooling proceeds, the representative point moves along the equilibrium
curve toward the point C. In a similar way, when we start from the
point N, the state of the cooled system first strikes the curve BB f and
then begins to move along it to the left. In no case, however, does
the representative point move beyond the intersection C of the two
curves. The point C is a triple point in which the liquid phase is in
equilibrium both with bismuth and with cadmium. As soon as it is
reached both components begin to freeze out simultaneously, forming
a microcrystalline mixture of the two solid phases (pure bismuth and
pure cadmium). The point C corresponds, therefore, to the lowest
188 TEXTBOOK OF THERMODYNAMICS X73
temperature at which the alloy can be maintained in a liquid state at
the given pressure p: for this reason, it is called the eutectic point
(eutectic = Greek for readily melting). The dotted parts of the
curves A A 1 and BB' have no physical reality as far as the system under
consideration is concerned. For comparison we give the experimental
Tjcdiagram of the cadmiumbismuth alloy 1 (solid curves in the same
Fig. 30) : considering the crude approximation of the formulas (10.09),
the agreement must be regarded as very good. The theoretical
eutectic point lies at the same concentration as the experimental
(x = 0.55) but about 25 too low.
This example is quite characteristic. The microcrystalline mixture
into which the liquid phase congeals at the eutectic point is called,
in the case of metallic components, the eutectic alloy. It has, usually,
the same content of the two components as the liquid, but this may
be said to be accidental, in the sense that it is due to the kinematics of
the freezing process. As we know from section 39, the equilibrium
conditions do not depend on the extension of the phases. Therefore,
the eutectic liquid should be in equilibrium with the solid phases taken
in any proportion. This is, in fact, observed and indicated in Fig. 30
by the horizontal line passing through the empirical eutectic point.
In the case of aqueous solutions of salts, the eutectic congelations were
formerly called cryohydrates.
It is worth pointing out that, as long as the solution can be regarded
as perfect, each of the two curves given by eqs. (10.12) depends on the
properties of one component only and is quite independent of the other.
In application to our example (CdBi) this means that the curve AA'
of cadmium has exactly the same shape and position, no matter what
the other component is. In order to measure a large part of it and to
bring out its curvature, we must choose, as the second component, a
metal with a low boiling point, e.g. mercury. Unfortunately the
7#diagram for CdHg was not available, and to illustrate this point
we give in Fig. 31 the diagram for zincmercury. When we consider
a mixture of vapors of two immiscible liquids in equilibrium with one
of the pure liquid phases, the corresponding curve gives us the equi
librium of this pure phase with its own vapor. The partial vapor
pressure of this component is, therefore, exactly the same as if the
other component were absent.
73. Case of partial mutual solubility. Other complications. In
order to complete our discussion of binary equilibrium, we shall give
here a brief schematic r6sum6 of the conditions which are to be expected
in the case of partial solubility in some of the phases. To fix our ideas
1 Petrenko and Fedorov, Zs. anorg. Chemie 6, p. 212, 1914.
X73 EQUILIBRIUM OF BINARY SYSTEMS 189
let us consider the case of two metals which form mixed crystals of
two kinds. The bulk of the kind a consists of the first metal while the
content of the second is smaller and can vary continuously from noth
ing to a certain limit. In the kind ft the conditions are reversed: the
second metal dominates and the proportion of the first has a continuous
range with a relatively small upper limit. Crystals of intermediate
composition do not exist. There are, therefore, two cases of solid
liquid equilibrium: coexistence of the liquid alloy of the two metals
with the mixed crystal a and with the mixed crystal ft.
As long as the mol numbers of the second metal (#2, #'2) are small
in both phases, they can change continuously and the conditions are
exactly the same, in this range, as in the case of complete solubility
(section 71). Assuming, by way of an approximation, that the solu
tion is perfect, we can describe the equilibrium between the liquid
alloy and the crystal a mathematically by the formulas (10.09),
(10.10), and (10.11). The graphical representation is, therefore, also
analogous to the corresponding range of Fig. 27. Two curves diverge
from the boiling point A of the pure first component (Fig. 32), the
one giving the mol number x\ in the liquid phase, the other x\ in the
crystal a; points of the two curves lying on the same horizontal are in
equilibrium. However, these curves do not extend across the whole
diagram but stop when x\ reaches its limiting value. This fact has a
bearing on the interpretation of the constants TOI, To2 of eqs. (10.9):
For the same reason as in section 71 we conclude that TOI is the
melting point of the pure first metal. However, we cannot identify
To2 with the melting point of the second because it is not permissible
to extrapolate our curves until this metal is pure. It is better to denote
this constant by r'o2, and to bear in mind that its numerical value
cannot be predicted from any general considerations. It can be larger
or smaller than TOI : When T'oz > TOI the two equilibrium curves go
from the point A upwards; when T'o2 < TOI, they go downwards.
Similar conditions prevail in the other case of equilibrium (the coex
istence of the liquid phase with the mixed crystal 0), when the mol
numbers x\, x'\ are small. The two equilibrium curves diverge from
the melting point B (temperature To2) of the pure second component
and are again represented by eqs. (10.9), (10.10), and (10.11) with
the new constants T'oi, To2 They go up or down from the point B
depending upon the numerical value of the ratio T'oi/To2 We have,
therefore, to distinguish three cases: (1) The equilibrium curves go
downward on both sides, from the point A and from the point B.
(2) They go downward on one side and upward on the other. (3) They
go upward on both sides.
190
TEXTBOOK OF THERMODYNAMICS
X73
The first case, illustrated by Fig. 32, is the most common. Suppose
we slowly cool the system, starting from a state represented by the
pair of points NN'. Some a crystal freezes out consisting mainly of
the first component; the remaining liquid therefore becomes richer in
the second component. The representative points move down along
the branches AD and AC, so that laterfreezing crystalline fractions
are of lower XL This continues until the eutectic point C is reached
marking the intersection between the " liquidus " curves AC and
BC. In this state the liquid alloy is simultaneously in equilibrium
with the mixed crystals a and ft. As in the case of the preceding
section, the alloy freezes completely when the temperature is lowered
further. The only difference is that it forms a microcrystalline mix
Liq.
FIG. 32. FIG. 33. FIG. 34.
Equilibrium in case of partial mutual solubility.
ture, not of the two pure substances, but of the two mixed crystals
a and 0, having the composition indicated by the points D and E.
This is the end of the motion along the branches AD, AC, and their
dotted continuation below the line DE has no physical reality. The
mixed crystals represented by the points D and E are not only in
equilibrium with the eutectic liquid alloy C but also with each other.
The question may be put, therefore, as to the equilibrium of the two
crystalline phases at temperatures below the eutetic. The answer is
given by the two schematic lines going downward from D and E:
Points of them lying on the same horizontals are in equilibrium. It is
an empirical fact that these lines slope away from each other.
The second case is represented by Fig. 33, which is selfexplanatory
because the notations in it are the same as in Fig. 32. We have used
the example of two metals only to fix our ideas. All we have said
applies, mutatis mutandis, to the equilibrium of two partially miscible
liquids with their vapor, and to similar cases. The cases illustrated by
the Figs. 32 and 33 are, in fact, as characteristic of mixed crystals as
of the binary equilibrium of liquids and vapors. However, the third
case represented by Fig. 34 applies only to a rather unusual system:
X74
EQUILIBRIUM OF BINARY SYSTEMS
191
Two substances form, in the solid phase, a mixed crystal in all pro
portions, but, upon being melted, they are partially miscible liquids.
At temperatures lower than the eutectic (C) the crystal melts either
into the solution a or the solution ft (depending upon its composition),
at the eutectic point into a turbid emulsion of the two solutions.
It goes without saying that the conditions considered here include
the limiting case when the first component is partially soluble in the
second but does not admit the second as a solute and crystallizes
(condenses or melts) in its pure state. The point D in the above figures
lies then on the vertical x\ = 1 passing through A.
We have enumerated in this chapter the types of binary equilibrium
which can be expected when the components do not form chemical
FIG. 35. Liquid us curve.
compounds. The presence of chemical compounds does not add much
new to the discussion, from the conceptual point of view, but makes
the curves look more complicated. One example will be sufficient to
illustrate this: In Fig. 35 we give the 7*diagram for the melting of
CaMg alloys. 1 These two metals form the compound CaaMgr,
there are three solid phases (pure Ca, pure Mg, CaaMg^ which happen
to be insoluble in one another. At 721 C we have the equilibrium of
the liquid and solid phase of the compound. The diagram is divided
by this point into two halves: The left representing a binary system
composed of the compound and magnesium, the right of the compound
and calcium.
74. Maxima and minima of the equilibrium curves. We have
seen in the preceding sections that the treatment of binary systems,
as if their phases were perfect solutions, gives a good qualitative
account of all the observed phenomena. It is even probable that the
not very large quantitative discrepancies are due, primarily, to the
crudity of our approximation in putting / = const (section 71) and
only in rare cases to the inadequacy of the description in terms of
1 1.C.T. baaed on Baar, Za. anorg. Chemie 70, p. 352, 1911.
192 TEXTBOOK OF THERMODYNAMICS X74
perfect solutions. In particular, we could account for the appearance
of maxima and minima in the equilibrium diagrams. However, it
will be useful to give here a theorem (due to Gibbs) relating to these
maxima and minima which is free from any assumption as to the
nature of the system.
Let us denote by <l>, Mi, M 2 the thermodynamical potential and
the masses of the two components in the first phase, by <', M'i, M f 2
the same quantities in the second. We shall use the equations of
equilibrium between the two phases in the form (6.33)
It was pointed out in sections 39 and 41 that $ is a homogeneous
function of the first degree in the mol numbers Ni, N% and the masses
Mi, M 2 . Consequently, the quotient </(Mi + M%) is a homogeneous
function of the degree zero; in other words, it depends only on the
ratio Mi/M2. In particular, we can represent it as a function f of the
mass fraction (10.05), and a similar conclusion holds with respect
to the other phase :
<*> = (Mi + M 2 ) f(yi), *' = (M'i + M' 2 ) f'(y'i), (10.14)
and by partial differentiation
 toi) + (1  *) ,  toO  yi . (10.15)
These expressions are quite general, no matter whether the compo
nents form chemical compounds or not. The variables entering into
eqs. (10.13) are, therefore, yi, y\, p, T. We give to these variables
increments dyi, dy'i, dp, dT in such a way that the two phases remain
in equilibrium, i.e. eqs. (10.13) continue to hold. We can, therefore,
take their total differentials
3A<f> ^
"^T 1 + ^7 ^ l + ~^T d + ~^F dT =  etc '
3^i 3^ i 3^> 3T
The partials with respect to p and T are given by the formulas
(6.51) and (6.52), while those with respect to yi, y'i follow from
eqs. (10.15)
X74 EQUILIBRIUM OF BINARY SYSTEMS 193
where AFi, AF 2 are the increases of the volume of the system and
l g i, I g 2 the latent heats contingent upon the respective transfers of one
gram of the first or second component from the unprimed to the
primed phase. We can eliminate dy'i by multiplying the first equation
by y ; i, the second by (1 y'i) and adding
Ox'i  yi) y^ d y* +A'dp B'dT = 0,
dy
i
with the abbreviations
A' = y'i&Vi + / 2 AF 2 , B f = (y'il g i + y'd&llT. (10.17)
We have now to distinguish two cases: (a) the ^diagram, at
constant temperature (dT = 0),
& y=iiL (10 .i8)
dyi A dyi*
(b) the 7>graph, at constant pressure (dp = 0),
dT y'i  yi
dyi B'
(10.19)
Obviously, the maximum or minimum of the py and !Pycurves
are at the places where the derivatives dp /dyi and dT/dyi vanish.
Since A', B' do not become infinite and d 2 /dyi 2 is not known ever
to be equal to zero, this condition reduces to
y'i = yi. (10.20)
The maxima and minima occur in the py and Tycurves at the same
concentrations, namely, when the mass fractions in the two phases
are equal. From the partial symmetry of eqs. (10.16) with respect to
yit y'\ it is clear that the conditions dp/dy'\ = and dT/dy r \ =
also reduce to the same eq. (10.20). This accounts for the fact that
the two curves (liquidus and solidus, or liquid and vapor) have always a
joint maximum or minimum.
Whether the condition (10.20) leads to a maximum or to a min
imum depends on the sign of the second derivative. The second
differentiation gives in the two cases (with y'\ y\ = 0)
d 2 p 1 d*$ d*T 1 d*f
^ dyi 2 A'dyi 2 ' x/ dyi 2 B' dyi 2
The quantities A' and B f can be considered as positive. In fact,
B' always is positive, since our notations are chosen so as to make the
194 TEXTBOOK OF THERMODYNAMICS X75
primed phase that of the higher heat a function (i.e. giving the positive
sign to the latent heats /i, fe). The changes of volume A7i, and AF 2
are always positive in the case of liquidvapor equilibrium, and posi
tive with rare exceptions in the case of solidliquid. The second
derivatives d?p/dyi 2 and d?T/dyi 2 have, therefore, generally opposite
signs. To a maximum in the pycurve there corresponds a minimum in
the Tycurve, and vice versa, as has already been mentioned in section 71.
It is well to point out that all these conclusions apply, independently
of the physical cause of the maximum or minimum type of curves, whether
they are due to association, compound formation, or other effects.
75. Remark on ternary systems. In a ternary system each phase
consists of three (independent) components Gi, 62, G$ with the mol
fractions #1, #2, #3 (or mass fractions yi, y^ ya) satisfying the condition
XI + X2 + *3  1. (10.21)
The scope of this book does not permit us to enter into the very
extensive theory of these systems. All we wish to do here is to explain
the type of diagrams used in connec
tion with them, in order to enable the
reader, to understand the graphical
material relating to ternary equi
librium. The triple of numbers xi,
X2, #3 can be interpreted as a point
within the area of an equilateral tri
angle. In fact, it is known from
geometry that the sum of the dis
tances of any internal point P from
the three sides (PAi + PA 2 + PA 3 ,
FIG. 36.Graphical representation in Fi *' 36 > is e( * ual tO the ^\ f
of the composition of ternary the triangle. If we choose the height
systems. as equal to unity, we may identify
the distances with the mass fractions
(PA\ = xi, PA2 = X2, PA$ = #3) because the relation (10.21) is
then satisfied. Every composition of the phase can, then, be repre
sented by a point of the triangle; and vice versa, every point within
it corresponds to a possible composition. The mol fractions x\ = PA\
of the component G\ are, usually, marked on the side GiG2, so that
one must draw the (dotted) line normal to PA\, in order to read yi
in the point 61. For instance in our Fig. 36, x\ has the value of
about 0.34. In a similar way the values of #2 (and #3) are listed on
the sides G*Gz (and GsGi) and can be read by drawing (dotted) lines
normal to PA* (and to PA*) to the points 62 (and 63).
X7S
EQUILIBRIUM OF BINARY SYSTEMS
195
When two ternary phases (primed and unprimed) are in equilibrium,
the condition (6.43) must be satisfied for each of the three components
Vk?k(p,T, xi, * 2 ) + 'k?'k(p,T, *'i, *' 2 ) = 0, (*  1, 2, 3)
because the mol fraction #3 (resp. #'3) should not be regarded as a
separate variable as it is completely determined by eq. (10.21).
Two of the variables, x'i, #' 2 or x\, # 2 , can be eliminated from
these three equations. If the pressure is, moreover, considered as con
stant, this leaves
T  r(*i, * 2 ) or T  r(*'i, *' 2 ), (10.22)
the equilibrium temperature expressed in terms of the mol numbers
of either phase. These equations depend on three variables (T, x\ t # 2
70,
70 60 50 40 30 20 10
FIG. 37. Isothermals of liquidus sur
face. (Temperature in C).
FIG. 38. Eutectic lines.
(Temperature in C).
or r, x'i, #'2) and can be represented graphically only in a three
dimensional space. For instance, we could interpret the third dimen
sion, normal to the plane of the triangle in Fig. 37, as the temperature
T. Equations (10.22) represent, then, two surfaces lying above the
triangle. If the equilibrium is between a solid and a liquid phase,
they are called, in analogy with the curves of Fig. 28, the soKdus
and the liquidus surface.
Naturally, one can represent in a plane drawing only the " levels "
of these surfaces, i.e. their intersections with the planes of constant
temperature, T = const. As an example we give in Fig. 37 1 the iso
thermal levels of the liquidus surface for the system NiFeMn (dotted
lines, the coordinates in this figure are not mol fractions but mass
fractions /i, y 2 , 3/3). In the solid state, these three metals are
1 1.C.T. based on Parravano, Gaz. Chim. Ital. 42 II, p. 367, 1912.
196 TEXTBOOK OF THERMODYNAMICS X75
mutually soluble in all proportions: therefore, there exists only one
solid phase with which the melt is in equilibrium and the liquidus
surface extends over the whole of the triangle. This is not the general
case: if the solid components are mutually insoluble or partially solu
ble, there may exist at the same temperature three (or even more)
different solid phases. This case is analogous to those of Figs. 30 and
32 for binary systems. As we have there two separate liquidus curves
corresponding to the equilibrium of the melt with the two solid phases,
so here the area of the triangle (Fig. 38) * is divided into three regions and
in each region we have a separate liquidus surface. The three liquidus
surfaces are analytically represented by three different equations
T = T'(y'\, y f 2) and each corresponds to the equilibrium of the melt
with one of the three solid phases. The lines of intersection of two
of these surfaces (eutectic lines) represent the states in which the melt
is in equilibrium with two solid phases simultaneously (Fig. 38). In
the point where all three intersect (ternary eutectic point) the melt is
in equilibrium with all three solid phases. There may arise further
complications in the number and shape of the equilibrium surfaces
when the components are able to form chemical compounds.
1 1.C.T. based on Goerens, Metallurgie 6, p. 537, 1909; Stead, Iron, Steel Inst. 91,
p. 140, 1915.
CHAPTER XI
FUGACITIES AND ACTIVITIES 1
76. Definition of fugacity. It was shown in section 42 that chemi
cal and physical equilibrium depends on the partial thermodynamic
potentials
We have reviewed, in Chapters VII to X, the cases in which some
theoretical knowledge of the functions permits the making of com
plete or partial predictions about the details of the equilibrium. How
ever, in the large majority of cases such knowledge is lacking or in
sufficient, and the only resource is to combine the experimental and
the theoretical methods: to measure and record the partial potentials
from one group of observations and to use these data for the prediction
of another set of phenomena. A large amount of experimental
material has been collected and published in the chemical literature.
However, the chemists usually record not the partial thermodynamic
potential itself but another function/ which stands to it in the simple
relation / = / (r) exp &/RT) (11.02)
Of V = RT[logf  log/ (r)]. (11.03)
The quantity / was introduced by G. N. Lewis 2 and is called the
fugacity or absolute activity. fo(T) is an auxiliary function of tem
perature only, which is defined by the following requirement: in the
gaseous phase, the fugacity approaches more and more the pressure,
as the density is decreased, and becomes identical with it (/ = p) in
the limit V = oo .
The thermodynamic potentials of perfect and of Van der Waals
gases were given by the formulas (5.41) and (5.44). In order
to satisfy the above requirement we have to put in both cases
f (T) =  w(T)/RT. For the perfect gas we obtain
/ = P, (11.04)
1 This chapter can be skipped without loss of continuity.
1 G. N. Lewis, Proc. Am. Acad. 37, p. 49, 1901; Zs. physik. Chemie, 38, p. 205,
1901.
197
198
TEXTBOOK OF THERMODYNAMICS
XI 76
at all temperatures; for the Van der Waals gas we find
2a RTb
(11.05)
Once/ is defined in the gaseous phase, it is definite also in the con
densed phases because (barring associations) the thermodynamic
potentials (and, therefore, the fugacities) are equal in equilibrium.
More generally, we have for two different conditions of the systems
(1) and (2), at the same temperature,
 /?  (11.06)
In general, analytical expressions for / are lacking and, even in
systems with one independent component, only the partial of log /
with respect to p can be given explicitly from (5.37)
(11.07)
The measurement of ^ can be easily carried out in pure
substances (systems with one independent component): in this case
s=^> = # Ts + pv, where each of the three terms can be directly
measured. As an example we give in Table 31 the fugacity of liquid
water at various temperatures and pressures. 1
TABLE 31
FUGACITY OF LIQUID WATER
Activity a
Fugacity /
25 C
37.5
50
25 C
37.5
50
1
1
1
1
0.03125
0.06372
0.1219
100
1.0757
1.0728
1.0703
0.03362
0.06836
0.13047
200
1.1576
1.1515
1.1461
0.03618
0.07337
0.13971
300
1.2454
.2356
1.2270
0.03892
0.07873
O.H957
400
1.3394
.3254
1.3132
0.04186
0.08445
0.16009
500
1.4402
.4214
1.4050
0.04501
0.09057
0.17127
600
1.5481
.5240
1.5029
0.04838
0.09711
0.18320
700
1.6637
.6336
1.6072
0.05199
0.10409
0. 19592
800
1.7874
1.7506
1.7184
0.05586
0.11155
0.20947
900
1.9200
1.8755
1.8367
0.06000
0.11951
0.22389
1000
2.0618
2.0089
1.9628
0.06443
0.12801
0.23927
1 M. Randall and B. Sosnick, J. Am. Chem. Soc. 50, p. 967, 1928.
XI 77 FUGACITIES AND ACTIVITIES 199
On the other hand, in a phase (i) with many components the
function <p h (f *, as is clear from its definition (5.39), is only then accessi
ble to direct measurement when the mol number JV* (i) can be changed
independently from the other mol numbers. If this is impossible
(owing to chemical reactions in the system), it is often still feasible
to determine it in a roundabout way when this component can be
isolated in another phase (j) which is in equilibrium with (i), making
use of the equilibrium condition ^ ( ) = ^ A ( ^. After being determined
from observations, these data can be used for theoretical purposes, for
instance to test the relations (6.49)
A*]ifctyk> =0, (11.08)
which must obtain within the phase (i).
77. Definition of activity. In some cases it is impossible or diffi
cult to determine the absolute value of the fugacity of a component
in a complex phase (for instance, in a solution), while it is feasible to
measure the ratio
*///. (11.09)
of the fugacities of the same component for two states of the solution,
different as to pressure and concentration, but of the same temperature.
This permits choosing /, as the unit in which the fugacity is to be
measured and regarding the state to which /. refers as the standard
state for the component and for the temperature in question. The
ratio a was also introduced by Lewis 1 and is called activity or relative
fugacity. Its relation to the partial thermodynamic potential follows
from the formula (11.03), taking into account that/and/, are measured
at the same temperature so that/o(r) drops out of the ratio:
?*?* = RTlo%a h , (11.10)
where v, h is the partial thermodynamic potential of the component h
in the standard state. The* simplest example is the case of a pure
substance: in Table 31 are given the activities of liquid water. As
standard states at the temperatures 25, 37.5, 50 C, are chosen
those corresponding to the pressure p = 1 atm (with/, equal to 0.03 125,
0.06372, 0.1219, respectively).
It must be distinctly understood that, in systems with several
components, the standard states are chosen for each component inde
pendently and arbitrarily. For the definition of ?*, not only
1 C. N. Lewis, Proc. Am. Acad. 43, p. 259, 1907; ZB. physik. Chemic 61 p. 129,
1907.
200 TEXTBOOK OF THERMODYNAMICS XI 77
but all the mol numbers of the other components, must be given.
Therefore, the <p,& are mutually exclusive, in the sense that there does
not exist any stable state of the system with these simultaneous values
of the partial potentials. This implies, of course, that they do not
satisfy the conditions (11.08). The expression
log K., (11.11)
1 h
does not vanish but is a finite function of temperature and pressure.
We find from (5.37)
AF,
' ajai ) &. <"'
/ p Ai
where
AF, =X) "A ^^* (? =]C vi^TldP* (1114)
On the other hand, the partial potentials^ in eqs. (11.10) refer to
actual states of equilibrium and satisfy the condition A$ = 0. There
fore, there follows the equation
]C ^ log a h u} = log X,, (11.15)
which has the same analytical form as the mass law, the mol fractions
being replaced by the activities.
The physical meaning of the activity becomes a little clearer from
the analogy which the formulas (11.10) and (11.15) bear to those valid
for dilute solutions. In section 59 we found for the partial thermo
dynamic potential of a dilute solute
Since log x* vanishes for x\ = 1, it may be said that the term ^
represents the molal thermodynamic potential of the fictitious state
when the component is pure while having the same properties \\& in the
solution. This means that the heat of solution is contained in <pn but
the entropy only partially: the second term RT log XH represents its
increase due to the entropy of mixing the solute (h) with the solvent.
The interactions of the component (h) with the other solutes and the
entropy of mixing it with them are negligible in dilute solutions. In
the general case of nondilute solutions, we have the analogous formula
(11.10)
>gaA 0) . (11.17)
XI 78 FUGACITIES AND ACTIVITIES 201
The standard state for the component h is, practically, always
chosen so that, in it, it is the only solute (xk = 0, k j& h). The term
with log a represents, therefore, the increment of the thermodynamic
potential due to interactions with the other solutes, to the entropy of
mixing h with them, and (in a measure depending on the definition of
the standard state) also to the interactions and entropy of mixing with
the solvent.
78. The activity coefficient. The simplest way of choosing the
standard potentials ,/ is identifying them with the functions <f>k of
eq. (11.16) used in the theory of dilute solutions
V*  ** w . (11.18)
The equilibrium constant K, becomes then identical with K in
Van t'Hoff s eq. (9.13), whence (11.15) takes the form
Yi "* a) log a h (i) logJC. (11.19)
J.h
The activities ah have then the meaning of effective mol numbers
which one must substitute instead of XH in order to conserve the form
of the mass law. The ratio between the effective and the actual mol
number
*h = Ok/Xk (11.20)
is called the activity coefficient. 1
In dilute solutions the activity coefficient is usually defined in
another way: as the ratio of activity and molality WH
y h = a h /m h . (11.21)
The relation of the two definitions (since in the dilute state
Xh = WA/WO) is
y h = CKA/WO, (11.22)
where mo is the molality of the solvent (i.e. mol number per 1000 g
weight).
It is obvious that the choice of the standard states defined by
eq. (11.18) is permissible when the solution obeys (in the limit of low
concentrations) the theory of dilute solutions treated in Chapter IX.
The standard states are then, really, possible states of the system,
and at low concentrations the activities approach the mol fractions:
lim ah = x ht lim ah 1. (11.23)
1 This term was first introduced by A. A. Noyes and W. C. Bray (Journ. Am.
Chem. Soc. 33, p. 1643, 1911), but these authors defined it with respect to the
molality as in eq. (11.21).
202 TEXTBOOK OF THERMODYNAMICS XI 78
However, even if this should not be so, if the values ,* = <PH cor
respond to no existing states of the system, they can still be used.
They are, then, merely auxiliary mathematical quantities with respect
to which the activities are defined. There is no reason why one
should not use thermodynamic potentials defined with respect to
fictitious states, provided there can be given an experimental way of
measuring them, either separately or in combinations. Such an experi
mental procedure exists and will be described below.
In the first place, it is simple to measure the activity of the solvent
by the method of boiling pressures or boiling (and freezing) temper
atures. For the process of vaporizing >o (2) mols of solvent, to obtain
FO (I) mols of vapor, we have from (11.19) the equation
(1)
vo log a (1)  vo (2) log ao (2)  log K, (11.24)
closely analogous to (9.17). We assume that the solutes in the liquid
phase (2) are nonvolatile so that the vapor phase (1) is pure and of
sufficiently low pressure to be regarded as a perfect gas. For a perfect
gas, eq. (11.16) or (9.09) is rigorous and the activity coincides with the
mol fraction: a* (1) = #A (I) = 1. This gives for the liquid phase (drop
ping the superscript)
vo log ao =  log K. (11.25)
In the particular case when the liquid phase also consists of the
pure solvent we can apply to it the theory of section 64 and find
log K(p,T) = 0, where p, T are the pressure and temperature of the
boiling point. When the solutes are present in a concentration that
is not high, the solution boils at a lower pressure p + &p (provided T
is the same), and we obtain from (11.25)
*o log ao   log K(p + A,!T)  AF A/> (11.26)
neglecting higher powers of A/>.
On the other hand, if the theory of section 64 were applied to
this case, we would have
j>o log XQ = AF
Denoting the ratio of the actually measured pressure increase A
to the theoretical Ao (if the solutes behaved like perfect gases) by
(11.27)
we have
log 00 = /3ok>g*o. (11.28)
XI 79 FUGACITIES AND ACTIVITIES 203
It is obvious, by analogy with section 64, that ft> can also be
measured as the ratio of the actual (#) and the theoretical (tfo) lower
ing of the freezing point, 0o = #/#() Much of the experimental work
has, in fact, been carried out by the freezing method. The above
formulas imply, of course, concentrations sufficiently low for higher
powers of A being neglected in the expansion (11.26). A more accu
rate formula for the activity of water as a solvent, taking in second and
third powers of the lowering # of the freezing point, is 1
logio ao   (421.00 + 0.1640 2  0.0037tf 3 ) X 10~ 5 . (11.29)
79. The activity function of electrolytes. Equations (11.28),
(11.29) give a convenient way of determining the activity of the sol
vent. As to the solutes, their activities can be measured directly if
they are volatile. This is, however, a comparatively rare case. A
very important field of investigation are the electrolytes whose ions
are mostly nonvolatile. It was shown by Lewis that valuable knowl
edge about their activities can be obtained from eq. (11.28) by means
of Duhem's relation (6.16). Let us consider the case that every
molecule of an electrolytic solute is completely dissociated into ions
according to the formula vid + . . . + v T G r G n = 0. If the mol
number of the neutral electrolyte dissolved was N n , those of the ions
are N* = VhN n , the sum being v e N n , (v 6 = v\ + . . . + v r ). Denoting
*  V '?\T = 1  *o, (1130)
(where NQ and XQ refer to the solvent), the mol fractions become
x k = VKX/V., (ft  1,2.. .r). (11.31)
In Duhem's equation (6.16) the partial thermodynamic potentials
<ph can be replaced by log a h because, in eq. (11.10) defining the activ
ity, the function ^. is independent of the mol fractions #*. As the
variable of differentiation we take x
3 log a h
Now, from (11.20) and (11.31): log a* = log a* + log x + const.
Substituting this for the activities of the ions,
Abel, Redlich, and Lengyel, Zs. phys. Chem. 132, p. 201, 1928.
204 TEXTBOOK OF THERMODYNAMICS XI 79
The expression
log 7  ~ vh log a* (11.32)
defines the quantity y called the activity function of the electrolyte
(sometimes also referred to as activity coefficient or activation function).
( i,) + i+,a. (n.33)
In the limit of zero concentration (x = 0) the solution is reduced
to the pure solvent (0o = 1) and it is experimentally ascertained that
lim log ah = lim log 7 = 0. The integration of (11.33) gives, there
fore,
log 7
_ /Ti + * * 9l g'
/o Lx jc 3^
In the particular case of considerable dilution, the expression
(11.28) may be used for log a , and, moreover, it is then permissible
to neglect terms of second order in the small quantities x and PQ.
Hence the limiting law for low concentrations
 go 3/3o
X
Since in dilute solutions x cc m (molality) this can be also written
The pioneer work of collecting data about activities of electrolytes
by the freezing method was largely carried out by American chemists
under the leadership of G. N. Lewis and A. A. Noyes. This material
confirmed the division of these substances in two groups discovered
earlier by other lines of study. 1
(A) Weak electrolytes (mostly organic acids and their salts) are
those which form aqueous solutions obeying, in their dilute state, the
theory of Chapter IX. This implies /So = 1, 7 = 1 for all low concen
trations, so that eq. (11.34) is identically satisfied (0 = 0).
(B) Strong electrolytes (the common strong acids and their salts)
which exhibit the electrolytic properties in their most typical form
1 There are two other methods of measuring the activities of electrolytes.
(1) Electrical measurements (conductivity or electromotive force), as described in
section 114. (2) Influence of other solutes on the solubility. This method is
limited to electrolytes of low solubility.
XI 80 FUGACITIES AND ACTIVITIES 205
do not obey the laws of dilute solutions. While 0o tends towards 1 in
the limit of infinite dilution (m = 0), it appreciably deviates from unity
even for very low molalities. In fact, the curve representing 0o = 0o(w)
has a vertical tangent at m = 0, i.e. QPo/dm = oo . Several empirical
formulas have been proposed for this function, among them
/So = 1  ^W H , (11.35)
which is supported theoretically (section 115) and represents the
experimental material as well as any of the others. This formula
permits one to carry out the integration of (11.34) and to find the
limiting law for the activity function
log 7 = km*. (11.36)
It will be better to postpone the discussion about the structure of
the factor k and about the agreement of this formula with the experi
mental data until section 115.
80. Activities in binary gas mixtures. The deviations from the
laws of perfect gases in electrolytes (preceding section) are due, pri
marily, to electric forces. It will be instructive to mention, as another
example, the deviations in gas mixtures which are caused by molecular
cohesion. The activities of a number of gases in mixtures have been
determined by several investigators from their chemical equilibrium
with liquid or solid substances, and a summary of their results is con
tained in a paper by Randall and Sosnick. 1 The activity coefficient
of a component is defined as a h = fh/fh (0) Xh where /* is the f ugacity of
the gas h in the mixture, /A (O) its fugacity in the pure state, at the
same pressure.
Randall and Sosnick confirm for binary gas mixtures an equation
which was found by Hildebrand 2 to hold for liquid solutions, namely,
log ai//i (0) *i) = (Bx 2 2 + Cx 2 * + Dxz* + . . .)/RT. (11.37)
The character of this function appears from the curves of Fig. 39 3
relating to mixtures of ethylene and argon at 25 C.
The only theoretical equation of state for gas mixtures is Lorentz'
extension of the Van der Waals formula given in section 5. Although
its validity is uncertain and, presumably, restricted to a narrow range,
it is interesting to note that it leads to an expression similar to (11.37).
Lorentz' equation is formally identical with the ordinary Van der
i M. Randall and B. Sosnick, J. Am. Chem. Soc. 50, p. 967, 1928.
'Hildebrand, Proc. Nat. Acad. Sci. 13, p. 267, 1927.
8 Randall and Sosnick, loc. cit.
206
TEXTBOOK OF THERMODYNAMICS
XI 80
Waals formula (1.21); only, the "molal volume" v has here a
different meaning, namely, v = V/N (where N = ffi + N*), while
the constants depend on the mol fractions as follows:
a
b
+ 2ai2XiX2 + (122X2*,
+ 2bi2XiX2 +
The thermodynamic potential is obtained by an obvious generaliza
tion of (5.44)
p +
W1 (D
f
c p idT T
0.1
Argon 0.1 0.2 0.3 0.4 0.5 0.6 0.7 0.8 0,9 Ethylene
Mol Fraction
FIG. 39. Logarithms of activity coefficients (log o&)
The entropy of the mixture must contain a term depending on the
mol numbers of both components which represents the entropy of
mixing the two gases. It can be evaluated by noticing that, in the
limit of very large v, the components become perfect. For a binary
mixture of perfect gases the thermodynamic potential is known to be
(section SO)
Ni*i(T) + N 2 <*2(T) + NRTlogp + RT(Ni
+ N 2 log * 2 ),
XI 80 FUGACITIES AND ACTIVITIES 207
where the subscript i stands for " ideal.*' Comparing the two expres
sions, we conclude that
(11.39)
Hence
It is our purpose to calculate the expression (11.37) to terms of
second order in the coefficients a, 6. Expanding and making use of
the equation of state (1.21), we find
According to (11.04), the fugacity of a perfect gas is equal to its
(partial) pressure, /u = pi = px\. Consequently (11.06) leads to
This relation holds for all values of the mol fractions, including
X2 = 0, xi = 1, when the left side is identical with log (/i (0) /) Tak
ing the difference of the two expressions, we obtain, in fact, an equation
of the type (11.37) with the following values of the coefficients.
B = p[(an  2ai2 + a 2 ^/RT  (6u  2b i2 + 622)]
 612) + an(a 2 2/2RT  6 22 ),]
\2
) ["
/
(3an  4ai2 + a 2 2)(ai2/2RT  612)
 & 22 )1 ,
3 . (an 
(611 2&i2 + 622)]
In the case of ethylene (1) and argon (2) at 25 C, the table values
are an/RT = 135, in 42, a 2 2/RT 53, & 2 2 30. The order of
magnitude of the data in Fig. 39 is consistent with this if a\2 and 612
208 TEXTBOOK OF THERMODYNAMICS XI 80
have commensurate values. At pressures of 25 atm the terms with f?
are still unimportant and, neglecting them, our formula is symmetrical
in the subscripts (1) and (2). This is borne out by the symmetrical
character of the curves for argon and ethylene at this pressure. The
constants can be adjusted so as to describe also the general character
of the curves at the higher pressures of 50 and 75 atm, but it remains
to be seen whether the expressions are fit for a quantitative descrip
tion of the phenomena.
CHAPTER XII
THE CAPILLARY LAYER 1
81. Surface layer and surface tension. The equilibrium of two
adjacent phases was treated in Chapter VI as if each of them were
uniform in its properties up to the very surface of discontinuity sep T
arating them from each other. This, however, is not strictly true:
the state of a substance at any point within it is influenced, not only
by the molecules nearest to this point, but also by all the others in a
small radius around it called the radius of molecular action. The sur
face layer of the thickness of this radius, between the phases or at the
boundary, is, therefore, in a condition different from interior points,
as it stands under the action of molecules of both media. It is called
the capillary layer and exhibits peculiar properties depending on the
nature of the two phases. When the phases are of large size, the
energy and entropy of the capillary layer become negligible, and the
conditions are those discussed in Chapter VI. But when the size of
the phases is small, in one or several dimensions, the influence of the
capillary layer is important.
A general characteristic of the capillary layer is that it tends to
contract. From the point of view of mechanics, it acts like a mem
brane in a state of uniform tension. By this we mean a condition of
the membrane (Fig. 40) in which every line element dl of its boundary
AB experiences a normal force F n = a dl to the inward, <r being the
constant tension. If we draw an imaginary line CD in the interior of
the membrane it is also under the action of two oppositely equal nor
mal forces of the magnitude a per unit length of the line. When the
membrane has a chance to contract, it performs work in doing so.
Suppose that the boundary line is displaced from the position AB to
A'B', and the element dl to dl r , through the distance AL, as a result of
the contraction. The work done against the element dl is, according
to (2.01), F n AL cos (w,L) = a dl&L cos (n,L), or it is equal to a multi
plied by the area which is swept over by dl in its displacement. Conse
quently, the total work is equal to <r multiplied by the area swept over
1 This chapter can be skipped without loss of continuity.
209
210 TEXTBOOK OF THERMODYNAMICS XII 81
by the whole boundary (which, in turn, equals the decrease of the
area 2 of the membrane) :
DW  <r<G, (12.01)
counting dS positive when it is an increase. It does not matter
whether the membrane is plane or curved.
\!\
KJI
B' 1 'B "2 (2)
FIG. 40. Work done by surface tension. FIG. 41. Capillary layer.
The tension <r of the capillary layer is called the surface tension.
Its numerical values for the more important substances are contained
in Table 32. A few additional data for solutions and for solids will
be found in sections 86 and 87.
The thickness of the capillary transition layer is not well defined
as it represents a continuous change of properties from one phase to
the other. In the interest of the next section we shall, therefore, pro
ceed as follows. We envisage a point in this layer and lay an imagi
nary geometrical surface AB through it and through all the other
points which are similarly situated in relation to the two adjacent
phases (Fig. 41). We lay a second imaginary surface, A\B\, parallel
to it, in the upper phase (1), at a distance d\ where that phase is
already uniform, and a third parallel surface, A^B^ in the lower
medium (2) where this medium is uniform. In all the variations
which we shall let the phases undergo, the distances di, dz from AB to
AiBi and ^282 shall be considered as constant and not subject to
variation. We shall restrict ourselves to the case when the distances
d are small compared with the radius of curvature of the layer.
Imagine now a wide l rigid and adiabatic envelope which contains
parts of both phases and cuts through the capillary layer. By our
device the space within it is divided into three regions, as labeled in
Fig. 41, and the total internal energy and entropy can be divided into
three parts pertaining to these three regions
U  U+ t/ (2 >+ /< 3 >, S = 5 (1 >+ S+ S< 3 >. (12.02)
1 Near the envelope the capillary layer is distorted because the walla add their
own surface effects; but when the envelope is sufficiently wide, the distorted parts
are negligible compared with the undistorted remainder.
XII 81
THE CAPILLARY LAYER
211
TABLE 32
SURFACE TENSIONS OF LIQUIDS IN CONTACT WITH GASES OR OTHER LIQUIDS
Substance
Adjacent medium
Temp
9 (dyne/cm)
Water
Air
C
75 49
Water
Water vapor
20
30
80
72.35
71.03
62.3
73 21
Water
Benzene
20
20
70.60
33
Water
Chloroform. . .
20
27 7
Water
Ethyl ether
20
9 7
Mercury
Air
20
471 6
Mercury
Vapor
100
360
20
456.2
376.4
480 3
Mercury
Ethyl ether
20
398 3
Mercury
Ethyl alcohol
20
364 3
Mercury
Chloroform .
20
356 6
Oxygen
Vapor
70 K
18 3
Nitrogen
Vapor .
90
70
13.2
10 53
Hydrogen
Vapor . .
90
15
6.16
2 83
Argon
Vapor
20
85
1.97
13 2
Carbon monoxide
Vapor
70
12 11
Carbon dioxide
Vapor
0C
9 13
Ethyl ether
Vapor
20
30
20
1.16
0.06
16 49
Benzene
Vapor
100
80
7.63
20 28
Methyl alcohol
Vapor
100
20
18.02
23 02
Acetic acid
Vapor
100
20
14.80
23 46
Carbon tetrachloride ....
Vapor
20
25 68
Ethyl alcohol
Vapor
100
20
16.48
23 03
100
14.67
The energy differentials in the first and second region (first phase
and second phase) have the usual expressions
(12.03)
212 TEXTBOOK OF THERMODYNAMICS XII 82
while that of the inhomogeneous intermediate layer is
rfZ7 ( 3 ) = TdSDW = TdS+ <rd2  &4V*\ (12.04)
where !T <3) and ps are the mean temperature and mean pressure in it.
82. Conditions of equilibrium as modified by the surface layer.
The fundamental conditions (5.03), in view of the relations (12.04),
take the form
0, 65 (1 >+ 6S+ 6S (3) = 0. (12.05)
When treating equilibrium with neglect of the capillary layer
(section 40) we found it convenient to divide the problem into two
parts by suitable restrictions of the virtual variations (auxiliary con
straints). We shall now go a step farther and break up the present
treatment into three partial problems.
(1) We do not admit any changes in the composition and mass or
in the volume and geometrical shape of the phases. The only per
missible variations are those of transfer of heat from one phase to
another. Under these conditions
6/o>= r<$S<, (j = 1,2,3).
We substitute this into the first eq. (12.05) and add the second,
multiplied by the Lagrangean factor X
o,
whence
jxn = r (2) = r< 3) = r =  x. (12.06)
In equilibrium the temperature is uniform throughout the system.
(2) We consider only variations at constant temperatures and
constant total volume. As was shown in section 36, the equilibrium
is then determined by the minimum of the sum of the work functions
whose differentials are (5.24)
Moreover, we disallow any changes of composition or mass in the
phases and permit only changes in volume and shape of the three
regions at the expense of one another. Since 67* = 0, the variation
d becomes, according to (12.03), (12.04)
v S  0.
XII 82
THE CAPILLARY LAYER
213
The variations of volume are subject to certain subsidiary condi
tions. In the first place the total volume is constant since we suppose
it enclosed in a rigid envelope
or
(pi 
 * 52 = 0.
(12.07)
As a further restriction, we specialize the geometrical shape of the
phases in the following manner : let the capillary layer have the con
stant curvature
C =  +
(12.08)
r and r r being its main radii, and let the walls of the vessel be normal to
the layer. We consider a transfer in which mass is added to the lower
phase (1) and taken from the upper
phase (2), so that all three surfaces
defining the layer are displaced up
ward through the distance /. If we
denote the respective areas of the
surfaces AB, AiBi, A 2 B 2 by 2, Si,
S 2 (Fig. 42)
6 7<D = 2i $/, 5 7 (2) =  S 2 81.
FIG. 42. Virtual displacement of
capillary layer.
On the other hand, we have from geometry
S/rr' = 2i/(r  di)(r'  di), etc,,
or neglecting squares of d/r, d/r',
2i = (1  Cdi) 2, 2 2  (1 + Cd 2 ) 2.
In the same way the increase of the area S due to the displacement
s
62 = CS6Z.
Substituting all this into (12.07), we find
pi  p2 = [<r  (ps ~ i)di  (#3  p2)d 2 ] C. (12.09)
The terms (3 p\)d\ and (3 pz)d 2 represent the difference
between the actual force of pressure on the unit length of the cross
section through the region (3), and the force which would act if the
phases (1) and (2) extended to the surface AB and changed discon
tinuously at it. Because of this subtraction, any indeterminateness,
214 TEXTBOOK OF THERMODYNAMICS XII 82
due to the arbitrary choice of the thickness di and </2, cancels out.
The difference itself is an intrinsic property of the capillary layer of
the same nature as the surface tension. In fact, there is no experiment
which would permit its measurement separately from it. It is, there
fore, appropriate to throw it together with the surface tension and to
denote by the symbol <r the whole bracket expression of (12.09). In
this sense, we write
p l  p 2 = Co = * +  (12.10)
It should be noted that in Fig. 42 the region (1) lies on the concave
side of the surface AB. We therefore make the rule that C is to be
counted as positive when (1) is on the concave side and as negative when
it is on the convex. Since <r is essentially positive, the pressure is
always higher on the concave side of the capillary layer. The relation
(12.10) is known as Lord Kelvin s formula because it was deduced by
that physicist from mechanical considerations (next section).
(3) We maintain constant the temperature T and the local pres
sures p\ 9 pz, p3 in all the three regions. As we know from section 36,
the equilibrium is then determined by the minimum of the thermo
dynamic potential : 6$ = 0. The only permissible variations are now
those of transfer of mass from one phase into another. Besides, we
restrict these variations still further by prohibiting any changes of
mass or composition in the region (3) and we consider only the trans
fers of mass of any component h (through the capillary layer) from
phase (1) into phase (2), or vice versa. These variations are precisely
the same as those considered in section 40, and the presence of the
capillary layer in no way influences the form of the conditions of
equilibrium. We may, therefore, refer to that section and to 42 and
take over the equations (6.43) and (6.45)
^"W^+^W^O, (12.11)
or
= ?W 2) . (12.12)
In the second of these equations the partial thermodynamic poten
tial is referred to 1 g of the component h (and not to 1 mol).
We may summarize the results of this section in the following way.
We derived the conditions of equilibrium for two phases separated by
a capillary layer. Comparing them with those obtained neglecting
the surface layer, we find all the conditions the same, except that
relating to the pressures of the phases. Instead of equality of pres
XII 83
THE CAPILLARY LAYER
21$
(1)
sure, we find a pressure difference (12.10) depending on the curvature
of the capillary layer. When the layer is plane, C = 0, its influence
on the equilibrium vanishes.
83. Direct observation of the capillary pressure difference. It
was known for a long time that the level of water is higher in narrow
tubes (capillaries) than in large vessels (Fig. 43). The theoretical
explanation of this phenomenon was given by Laplace, who showed
that the column of liquid in the capillary tube is supported by the
surface layer of the meniscus. In this way the names " capillary
layer " and " capillarity " originated. A particularly simple vffcy of
calculating the difference / of levels in the tube and
in the wide vessel (" capillary rise ") was given by
Lord Kelvin. Let us consider the case that a
liquid rises in a tube of circular crosssection
(with the radius R) and that the meniscus includes
the angle # with the walls. The vertical com
ponent of the force of tension which acts on the
rim of the capillary layer is then F = 2irR<r cos $.
This force supports the liquid column of height /
and must be equal to its weight in the surround
ing gaseous atmosphere. Denoting the densities
of the liquid and gas by P2 and pi, respectively,
and by g the acceleration of gravity, F = irR 2 l g(p% pi), whence
2<rcos# = Rlg(p2 PI). For simplicity, let us assume the curvature
1(2)
FIG. 43. Capillary
rise.
of the meniscus as constant over its whole surface,
curvature is then r = R/cos &, and we find
The radius of
 Pi)
(12.13)
For water and other wetting liquids # = and, roughly, R = r .
On the other hand, for completely nonwetting liquids, like mercury,
& = IT: i.e. r and / are negative. This means that the liquid is on the
concave side of the meniscus and its level is lower in the capillary.
The formula (12.13) offers one of the most accurate methods of measur
ing surface tensions.
If p is the pressure at the plane surface in the vessel, the pressure
in the gas at the meniscus is smaller by the pressure of the layer of
gas of the thickness I, viz. pig/:
Pi
(12.14)
216 TEXTBOOK OF THERMODYNAMICS XII 84
Similarly the pressure pz in the liquid at the meniscus is decreased
. . 2<r P2
p 2  p =
r pi p2
Consequently
Pi p2= (12.15)
This is identical with the thermodynamical formula (12.10), in the
special case r = r'.
84. Influence of curvature on the pressure and temperature of
transformations. Let the pressure and temperature of equilibrium
between two phases be p and T when the capillary layer is flat (wide
vessel). We first consider the case of a pure substance (or of a system
with one independent component), <p = (p, and use the equation of
equilibrium in the form (12.12), since we are not interested in the
effects of association, p and T must then satisfy the relation
(12.16)
When the capillary layer is curved, the conditions are slightly dif
ferent. We denote the temperature T + Ar and we have two pres
sures pi = p + &p and p 2 = p + A/> <rC, according to (12.10).
The relation of equilibrium still has the same form
<P M (P + A/>  <rC, r + Ar)  9M ^(p + A/>, r + Ar) = o.
Restricting ourselves to quantities of the first order in the small
increments, we expand into a Taylor series and subtract (12.16)
This formula can be specialized in two ways:
(1) Transformation pressure at given temperature. Suppose that
the flat and the curved boundaries are at the same temperature:
Ar = 0, so that the third term of the equation vanishes. According
to section 35, (9vW9/>)T ** ^M, the specific volume referred to Ig, which
is the reciprocal of the density, V M = 1/p. Hence
Pl (12.17)
'
PI ~ P2
In the special case of equilibrium between a liquid and its vapor
this is identical with the relation (12.14) which was obtained by Lord
XII 84 THE CAPILLARY LAYER 217
Kelvin's method. In fact, the liquid boils at the meniscus in the tube
at a slightly different pressure from that at the plane surface of the
vessel, corresponding to the difference of altitudes.
(2) Transformation temperature at given pressure. Suppose the
pressure of the phase (1) is the same at its plane and at its curved
boundaries: pi = p or Ap = 0. We know from section 42 that
\ /W>\ _ I
),~ \~&r) f  ~
\~&r,~ \~&
where IM is the latent heat of transformation at the plane surface, i.e.
the heat necessary to transform Ig of the substance from the phase (2)
into the phase (1),
AT = ~ (12.19)
P2/AT
In addition to the case of a pure substance, we shall be interested
in another which is but slightly more general. We assume that one
of the phases, say (2), is condensed and pure, containing only the com
ponent h, while the other phase (1) is either a gas mixture or a dilute
solution. We consider first the case of the gas mixture and apply to
it eq. (9.24), since it can be considered an immediate consequence of
(6.43) or (12.11),
RT log p, = fok^tor)/**  *(r)J, (12.20)
where g* = ^ (1) /^ (2) . At a flat surface of discontinuity, p means the
pressure in phase (2) identical with that in phase (1) where it is com
posed of the partial pressure ph (l) of the component (h) and of the pres
sures of all the other components whose sum we shall denote by po, viz.
p = p h w f pQ. The curvature of the surface of discontinuity cannot
influence the pressure p^of those components which are absent in the
pure condensed phase. It remains the same, and only the partial
pressure of the component h changes into ph (l} + A/> A (1) . Consequently
p 1 = p + A/> A (1) and p 2 = p +' A/> A (1)  aC, according to (12.10). At
the temperature T, we have, therefore, from (12.20)
> aC, T)/g h 
or, as above,
When the first phase is a gas mixture (ph (1> = RT/vS 1 *), this is iden
tical with eq. (12.14) since obviously (&ifc (1) Vh^/Vk* = (P2 PI)/DI.
218 TEXTBOOK OF THERMODYNAMICS XII 85
In the case of a dilute solution, (1) labels the liquid phase and (2) the
solid precipitate. ph (1) must be interpreted as the osmotic pressure of
the component h which is proportional to the solubility x gh a \ according
to the relation (9.50). Neglecting A/>* in the parentheses beside aC,
(where C is negative, since the solid particles are convex)
,
which expresses the influence of the curvature upon the solubility and
is called the OstwaldFreundlich formula. It is to be remembered that
v h (2) is the molal volume in the condensed phase. If the substance is
dissolved without change of molecule (g h = 1), the solubilities for the
two negative curvatures C\ and 2 are related by the formula
log*.u (1)  log*.A2 (1) =  ^ (2 V(Ci  C 2 )/RT. (12.22)
8S. Applications. The effects listed in the preceding section have
an appreciable magnitude only in the case of very large curvatures.
Let us take as an example drops of water at room temperature. The
constants for water at 20 C are v M (2) = 1, jr (1) 5.0 X 10 4 cm 3 /g,
1 M = 2.45 X 10 l erg/g, <r = 72.3 dyne/cm. Thence and from (12.19),
with C =  2/r,
(12 .23)
Even for tiny droplets of the radius r = 10 ~ 5 cm the boiling point
is only 0.174 degree below that at a plane surface. However, the sur
face layer is only two or three molecular layers thick, so that our for
mula is quantitative down to about r = 5 X 10~ 7 cm.
Why then pay much attention to this apparently insignificant phe
nomenon? Because it plays a large role in the processes of trans
formation of state, as they actually occur in nature. Let us consider,
for instance, the condensation of water vapor when its temperature
falls slightly below the point of saturation. Granting that the boun
dary layers of the vapor can and do condense on the surface of the
container, its main mass is spread through the volume. Here the
condensation must first form a liquid nucleus before it can proceed
further, a tiny droplet which has a low boiling point, according to
eq. (12.23), and is, therefore, unstable. Instead of growing it is more
likely to dissolve by vaporization. This is the reason why condensa
tion does not take place through the volume of a completely pure
vapor. Only when dust particles are present, do they offer to the
vapor a surface of reasonably low curvature and act as nuclei of con
XII 85 THE CAPILLARY LAYER 219
densation. Very effective in this respect are electrically charged
particles: their field produces an electrostatic tension at the surface of
the condensed droplet which is inversely proportional to the fourth
power of its radius 1 . When the radius is very small, it may overcom
pensate the effect of the curvature and make the condensation easier
than at the plane surface.
Quite similar conditions obtain in the case of vaporization. When
water boils at 100 C, the outer pressure does not prevent small vapor
bubbles arising in it. However, from the point of view of the water,
the surface of the bubbles has a negative curvature, so that the boiling
point in them is above 100 and they are unstable. Only when bub
bles of gases already exist in the water from the start, does the boiling
take place through its volume.
In the same way, surface tension has an effect on transformations
in condensed phases. It is well known that crystallization does not
begin without suitable nuclei, and the explanation is the same as for
the condensation of gases. The case of fusion is also interesting, since
most solids consist of very small microcrystals. The phenomena in
crystals are somewhat complicated by the fact that the different
crystallographic planes have different surface tensions, but, qualita
tively, they are similar to those treated above. We should, therefore,
expect that, with the rise of temperature, the smallest units and the
sharp corners and edges of the larger ones would melt first (because
these parts have the lowest point of fusion) while the larger crystals
swim loosely in this melt. This influence of microcrystalline structure
upon the process of fusion has not yet been sufficiently appreciated.
However, it seems to have a bearing on the observation of Errera that
glacial acetic acid and a few other organic solids show an abnormally
high dielectric constant two or three degrees below their melting
points. 2
1 The electrostatic tension or additional difference in pressure between the atmos
phere and the interior of the droplet, due to the charge e on it, is equal to e*/Sira 4 .
This term must be added to aC in eq. (12.17).
2 J. Errera (Trans. Far. Soc. 24, p. 162, 1928) determined the dispersion curve
of the dielectric constant in the vicinity of the melting point and found that the effect
in acetic acid has its maximum at 3 below 7> and for a critical frequency of electro
magnetic waves v = 10 3 . Assuming that the phenomenon is due to free solid particles
capable of rotating in the liquid melt, my colleague Prof. G. W. Potapenko calculated
from the critical frequency the mean radius of the particles and obtained r  3 X 10 ~ 8
cm. This result agrees well with the formula (12.19) which gives for the surface
tension <r = Jrp 2 /jif &T/T. Substituting A r = 3 (as found by Errera), IM 1.9
X 10* erg/g, P2  1.266 we obtain (at T  298) a  40 dyne/cm as the surface
tension of solid acetic acid. This value lies in the range found experimentally and
theoretically for other solids (section 86),
220 TEXTBOOK OF THERMODYNAMICS XII 86
Attempts to use eq. (12.22) as a means for determining the
surface tension of solid salts against their saturated solutions were not
very successful. The solubility of barium sulfate and calcium sulfate,
in several states of dispersion, was measured by determining the elec
tric conductivity of the solutions 1 which increased with the dispersion.
But the very high surface tensions calculated from these data 2 have
been questioned since there are possible sources of error which would
produce a similar rise of the conductivity. 3
86. Formulas for surface tension. Only when the molecular
forces holding a substance together are completely known, is the
theoretical calculation of its surface tension possible. So far, this
knowledge is available only in the case of crystalline solids and, even
here, it is restricted to a few simple space lattices. For heteropolar
halogen compounds of the type of NaCl (rock salt), KC1 (sylvine), etc.
Born and Stern 4 calculated the following expression
er = 4.022 X 10 3 . * . , (12.24)
Ai + A 2
where p is the density of the crystal and A\ t A% the atomic weights.
This gives:
TABLE 33
Substance LiCl NaCl KC1 NaBr KI
<r (dyne/cm)... 320 149 108 125 76
Although the BornStern formula is only an approximation, the
values derived from it are, presumably, as accurate as the results of
the very difficult experiments with surface tensions of plastic solids.
Experiments of this kind gave for pitch, at 20 C, about 50 dyne/cm
(Ignatiew), and for lead glass, at 500 C, 70 dyne/cm (Berggren).
Eotvos 5 proposed a formula for the tension of liquids against their
own vapor which he obtained by a partially theoretical reasoning. He
starts from the assumption that the extended law of corresponding
states of section 27 applies also to the surface energy <r S which must
be treated on the same footing as the volume energy pV or the internal
energy U. In other words, for substances which obey the law of cor
1 Hulett, Zs. phys. Chcm. 37, p. 385, 1901; Dunadon and Mack, J. Am. Chem.
Soc. 45, p. 2479, 1923.
W. J. Jones, Zs. phys. Chem. 82, p. 448, 1913.
Balarew, Zs. anorgan. Chem. 154, p. 170, 1926.
* M. Born, Verh. Deutsch. phys. Ges. 21, pp. 13, 533, 1919,
8 R. Edtvos, Ann. Phys. 27, p. 448, 1886.
XII 86
THE CAPILLARY LAYER
221
responding states, the ratio cr 2/r must be a universal function of the
reduced variables TT, T. If we imagine 1 mol of the liquid substance
expanding with the temperature in such a way as to retain its geometri
cal similitude, its free surface S will change proportionally to v*.
Therefore, the extended law of corresponding states leads to
av** _
~r~ ^ ' (7r ' T '*
When the function / is found for one substance, it is the same for
all others obeying the law of correspondence. From experiments with
ethyl ether between and 190 C, Eotvos inferred that the function
is/(7r, r) = 2.22(l/r  1) = 2.22(r c /r  1). Hence
<rv" = 2.22 (T e  T)
(12.25)
must be a universal relation between surface energy and temperature.
In fact, this formula, due to Eotvos, holds with a fair approximation
for numerous substances.
Ramsey and Shields 1 secured a better fit with experiments by
writing
ov* = k(T c  T  d), (12.26)
where k is slightly variable (but not greatly different from Eotvos'
value) and d is a new constant (Table 34). The formula loses, of
course, its theoretical significance and becomes empirical.
TABLE 34
Substance
t c C
k
d
Range in C
Ethyl ether
194.5
2.1716
8 5
20160
Methyl formate
214
2.0419
5 9
20190
Ethyl acetate
251
2.2256
6 7
20200
Carbon tetrachloride
283
2 1052
6
80250
Benzene
288.5
2.1043
6 5
80250
Chlorobenzene
360
2 077
6 3
150300
A purely empirical formula, giving the surface tension as a function
of temperature, is due to Van der Waals 2
(12.27)
1 W. Ramsey and J. Shields, Zs. phys. Chem. 12, p. 433, 1893.
2 J. D. Van der Waals, Proc. Amsterdam, 1893.
222 TEXTBOOK OF THERMODYNAMICS XII 87
It has been extensively tested by Verschaffelt l and represents the
data for a number of substances fairly well.
TABLE 35
Ether logio A = 1.761, B = 1.270
Benzene 1.839, 1.230
Chlorobenzene 1.810, 1.214
Carbon tetrachloride 1 . 445, 1 . 185
87. Influence of temperature and of impurities. The experimen
tal data on surface tensions in solutions can be summarized as follows.
All the solutes can be divided into two classes:
(A) Inactive solutes produce a slight increase of o, proportional to
ttieir molality nth (provided the concentration is not high)
ff = <r (l + km*), (12.28)
where ox> is the surface tension of the pure solvent. Alkali salts are
inactive in aqueous solutions and the coefficient k is of the same order
of magnitude for all of them, namely, between 0.0174 and 0.0357.
Similar conditions obtain for solutions of aromatic substances, like
camphor, aniline, benzoic acid in organic solvents.
(B) Active solutes have the opposite property of strongly lowering
the surface tension. Such are many organic substances in aqueous
solutions, for instance, alcohols, aldehydes, ethers, fatty acids, ter
penes, etc. Soaps are particularly effective: a solution of sodium
oleate of the molality m^ = 0.002 has a surface tension <r = 25 dyne/cm,
while that of pure water is <TO = 73. The cleaning properties of soap
solutions rest on the low values of their surface tensions.
In a qualitative way these facts were explained already by Gibbs.
Let us write down the expression for the differential of internal energy
of the surface layer. We consider the case that the mass of the system
is variable and write dU in a form analogous to (5.06), taking into
account (12.01) and (5.40)
dt/< 3) = TdS+ <r dS + ?*W 3 >. (12.29)
ft
The term pdV does not appear in this expression because it is
included in 0dS (as far as the surface layer is concerned), as was
explained in discussing the formula (12.09). Similarly, N h represents
here the difference between the actual value which the mol number of
the components h has in the whole system and that which it would
1 J. E. Verschaffelt, Mitteil. Naturk. Labor. Gent No. 2, pp. 16, 19, 1925.
XII 87 THE CAPILLARY LAYER 223
have if the surface layer were replaced by a sharp discontinuity. In
other words, Nh (3) is the quantity of the component h (in mols) adsorbed
by the surface layer. Although not expressly stated in section 82, it
follows from it that ^ in a surface layer of negligible curvature is the
same as in the two adjacent volumes. For the further discussion it is
convenient to introduce the generalized work function
whose differential is
^(3) =  sdT  W 3) <ty k + <* <&. (12.31)
h
We now make the assumption (borne out by all observations) that
the work function has the analytical form ^ (3) = fa S, where fa, the
work function per unit area, is a specific thermodynamical characteristic
of the surface layer independent of its area S. Substituting into
(12.31), we obtain
o = fa. (12.32)
The surface tension is identical with the specific work function per
unit area of the capillary layer. Moreover, we find
We apply the last formula to the special case when one of the
phases, adjacent to the surface layer and containing the component
(h), is either a perfect gas or a dilute solution in the sense of Chap
ter IX. According to (9.09) and (9.10), <p h = <p h + RT\ogx h , where
<ph is a function of p and T only. The partial with respect to 53* in
(12.33) is taken in the sense that all the other variables (including
T and p) are to be kept constant: 3^ = RT'd log #&. In dilute solu
tions Xh can be replaced by WA/WO, so that
This formula is the key to the behavior of solutes. (A) In the
case of inactive solutes: fiff/dm*, > 0, whence n^ h < 0, i.e. the adsorp
tion (per unit area) of the capillary layer is negative. This means
that the concentration of the component (h) is less in the capillary
layer than in the bulk of the solution. As we deal here with dilute
solutions, the solute can exercise but a small effect upon the surface
tension, in agreement with the observed facts. (B) In the case of
224 TEXTBOOK OF THERMODYNAMICS XII 87
active solutes there follows n^h > 0, the adsorption is positive. The
surface layer is very much smaller in volume than the adjacent phases
and may build up to a high concentration in the solute (A), by adsorb
ing it even from a very dilute solution. This is the explanation why
the increase of the surface tension by a small amount of solute is
always slight, while the decrease may be considerable.
It will be well to say here a few words about the temperature
dependence of the surface tension. As the data of Table 32 indicate, <r
decreases with rising T. Moreover, it may be concluded from the
good fit of the formulas (12.25) and (12.26) that this drop is linear
over wide ranges of temperature. In fact, the molal heat v appearing
in these expressions changes but little with T and can be regarded as
constant in the first approximation. The underlying cause of the
linear drop can be inferred from the following considerations. Let us
apply eq. (12.30) to a pure substance: the question of adsorption has,
then, no interest and we may leave out the last term, writing it in the
form (dividing by S) fa = w 2  7s s or from (12.32), (12.33)
Within the range where a can be represented as a linear function
(<r = cri aT) this gives
u% = <7i = const.
The internal energy (per unit area) of the capillary layer is constant
and independent of temperature. This ceases to be true in the vicinity
of the critical temperature T where the linear laws (12.25) and (12.26)
are no longer valid.
The understanding of the physical nature of certain surface films
(oil films) has been considerably advanced by investigations of Harkins
(and collaborators), Langmuir, and Adam, 1 but this subject lies out
side the scope of the present book.
1 W. D. Harkins, F. H. Brown, and E. C. H. Davies, J. Am. Chem. Soc. 39,
p. 354, 1917; I. Langmuir, ibidem 39, p. 1848, 1917; N. K. Adam, Proc. Roy. Soc.
(A) 99, p. 336, 1921; 101, pp. 452, 516, 1922; 103, pp. 676, 687, 1923.
CHAPTER XIII
NERNST'S HEAT POSTULATE OR THE THIRD LAW
OF THERMODYNAMICS
88. The principle of Thomsen and Berthelot. It was emphasized
in Chapter IV that the fundamental eq. (4.07) by means of which the
concept of the entropy is derived from the second law is a differential
equation. Because of this, the definition (4.09) of S contains an
additive constant of integration so that the absolute value of the entropy
cannot be determined from the first and second law. The indeter
minateness of the entropy constant has serious implications for the
general character of the thermodynamical laws, in that they contain
coefficients of indeterminate numerical value. As an example, we can
point to the law (8.17), (8.16) of chemical equilibrium of perfect gases:
The expression of the equilibrium constant K contains a factor /,
which depends, according to the formula (8.14), upon the entropy
constants of the gases taking part in the reaction. It is clear from
this that the knowledge of the absolute entropy would be of inesti
mable value, as it would greatly enhance the usefulness of many
thermodynamical formulas. To obtain this knowledge one must
have recourse either to experiment, or to new theoretical principles
not contained in the first and second law. Of course, it is quite suf
ficient to determine the entropy of a substance in one special case, in
order to deduce the entropy constant, and it was pointed out by
Nernst (1906) that experimental data were available to make a
general conclusion about the behavior of the entropy of condensed
systems in the vicinity of the absolute zero point of temperature.
Let us start our considerations from the GibbsHelmholtz equa
tion (5.48) relating to an isothermal process
We have seen in section 36 that, under ordinary conditions, the
process (reaction) takes place spontaneously when A$>0. The dif
ference of thermodynamic potentials A* m $ 2 $1 can be regarded,
therefore, as a measure of the chemical affinity. Suppose now that
226 TEXTBOOK OF THERMODYNAMICS XIII 89
the partial (d&$/dT) p does not become infinite; when T goes to zero.
In this case we find, putting T = 0,
AX) = 0, (13.02)
or lim Tm0 (&$ Q p ) = 0, if the process is conducted also at constant
pressure.
In the vicinity of the absolute zero point the heat of reaction
becomes identical with A< and the criterion of the reaction taking
place can be replaced by AX > or Q p > 0. For a long time it
was thought that this form of the criterion holds good at any tem
perature. In fact, Thomsen and, independently, Berthelot stated the
principle that out of a number of possible reactions the one will take place
which evolves the largest amount of heat. According to our present
knowledge, this would be true if the right side of eq. (13.01) were
always negligible so that we could write at all temperatures
A$  AX = 0. (13.03)
As we have already mentioned in section 37, this equation is
erroneous. Modern investigations have shown that the term
T(d&$/dT) p must be taken into consideration. Nevertheless, the
approximation given by the principle of ThomsenBerthelot, and by
eq. (13.03), is surprisingly good even at room temperature, provided
the system is condensed and the affinity of the reaction not very small.
A comprehensive set of measurements of A<i> = $2 3>i for reactions
in condensed systems was carried out by T. W. Richards 1 with the
result that the difference A< AX was very small, indeed.
Exercise 99. Check that at room temperatures r(DAf>/3D p is small compared
with A* for the reactions in galvanic cells given in formula (5.52) and in exercises
56, 57.
89. Nernst's own formulation of his postulate. This last fact led
Nernst to assume that in condensed systems the right side of eq. (13.01)
decreases more rapidly, when the temperature goes to zero, than is
accounted for by the factor T. According to him, the remaining
factor also decreases to naught in the limit T = : for any isothermal
process
The assumption expressed in this formula is known as Nernst's
postulate or the third law of thermodynamics. That it is natural and
* T. W. Richards, Zs. phys. Chemie 38, p. 293, 1902,
XIII 89 NERNSTS HEAT POSTULATE 227
convincing is best brought out by the following considerations. Equa
tion (13.01) can be written in the following form
If the partial on the left side does not become infinite, there follows
from (13.01) that both AS> AX and T become zero in the limit
!T = 0. It must be well understood what a partial with respect to T
at constant pressure means in this case. The pressure in the two terms
of A$ = $2 $1 and AX = X2 Xi is not necessarily the same:
A<i> = <S>2(r, 2 )  3>i(r,i), etc. But the partial is taken so that it
remains equal to p% in the first term and to p\ in the second. Let us
now take the limiting values of both sides of eq. (13.05) for T 0,
keeping the pressure constant (i.e. equal to p% in <i>2 and X2 and to
pi in <J>i and Xi). The right side is indeterminate, being equal to 0/0,
but the indeterminacy can be readily resolved by means of the well
known mathematical rule : instead of the numerator and the denomina
tor, one has to substitute, respectively, their partials with respect to T.
/9A$\ /3AX
mr  w hmr  Var /," hmr 
whence
0. (13.06)
If we plot the change of the heat function AX in any isothermal,
chemical or physical, process against the temperature T at which it is
conducted, the curve has a horizontal tangent at
the limit T = (Fig. 44). We repeat that this
conclusion is independent of Nernsfs postulate'.
the only assumption on which it rests is that
OAS/Q^Op does not become infinite for T = 0.
The difference A<J> is equal to AX in the limit
T = 0, and is very little different from AX over
a considerable range of temperatures, according FlG 44.
to the experimental data mentioned in the preced postulate.
ing section. Nernst's heat postulate (13.04) has
the purpose of insuring such behavior. The curves for A< and for
AX have, if it is accepted, a common horizontal tangent, at T = 0,
and remain for some distance close to each other.
Owing to eq. (5.37), (d$/dT) p S, an alternative form of
Nernst's postulate (13.04) is
lim r .o(5 2  Si)  0. (13.07)
228 TEXTBOOK OF THERMODYNAMICS XIII 90
The entropy remains unchanged in any isothermal process taking place
in a condensed system in the vicinity of the absolute zero.
Although Nernst enunciated his postulate only for condensed
systems, this restriction may now be dropped for the following reasons:
(1) In sufficient proximity to the zero point of temperature every known
system of nature is condensed, if the pressure is finite. (2) We shall
see in Chapter XVI that even the ideal equation of perfect gases has
been changed by the modern theory of statistics so as to satisfy
Nernst's law.
Exercise 100. Starting from eq. (5.58) prove:
(a) If limroteA^/aTV ^ , then lim^.o (&&U/dT)v  0.
Use the same argument as that used in the text for proving (13.06).
(b) From (a) and from (13.06) there follows
limro (CV2  CVO  0, Hmy.o (C P2  C p i) = 0.
These results are independent of Nernst's postulate.
(c) Equation (13.07) requires
0.
90. Thermal expansion in the vicinity of the absolute zero. There
can be deduced from the formula (13.07) some immediate consequences
about the properties which all substances have in their condensed state
at temperatures close to T = 0. Let the process to which we apply
this equation consist, merely, in an infinitesimal increase of pres
sure (from pi = p to p2 = p + dp) at constant temperature. The
entropy of the system is changed, accordingly, from Si = S to
S 2 = S + (dS/dp) T dp, and Nernst's postulate (13.07) takes the form
lim r .oO5/3/>)r = 0. (13.08)
Because of the relation (4.46) of section 28 this is identical with
lim r . () =0. (13.09)
The coefficient of thermal expansion vanishes in the vicinity of T = 0.
This fact is important in itself and leads to some interesting inferences.
If the pressure is kept constant, the volume is sure to remain constant
automatically: in the limit T = 0, a partial with respect to T at
constant p is, at the same time, a partial at constant v. In the formulas
of the last sections we could replace the subscript p by v or drop it
altogether. According to the definition (3.08) of the specific heat
this applies also to the quantities c p ,c v ,
lim r _ C P =* lini r .o d Hm r . c. (13.10)
XIII 91 NERNST'S HEAT POSTULATE 229
We can readily obtain the generalization of the formula (13.08)
for the case when the system is not simple but is described by the
variables Xi, J^2, X n and the generalized forces y\, 3^2, ... ^n, in
addition to T (compare section 7). In fact, let us consider processes
which consist in changing only one parameter, while the rest of them
remain constant: for instance, Xh changes to Xh + dX^ while all the
other Xj (j j h) and all the yj (including yh) stay constant; or else,
y h changes to yh + dyn and y, (j 7* h) and all the Xj are constant. We
find as above, respectively,
oC oC
lim r .o ^T  O t lim r . f  0. (13.11)
dXh Qyh
Because of (5.30) and (5.45) this is equivalent to
Hm,.o  0. fonro ?  0. (13.12)
The temperature gradients of all the parameters X^ yh describing the
em vanish in the vicinity of T = 0. In particular, we conclude that
system
91. Specific heat in the vicinity of the absolute zero. Nernst's
postulate implies that the entropy is finite at all (finite) temperatures,
including the limiting case T = 0. We can choose, therefore, T =
as the lower limit of the integrals (4.33), (4.34) representing the
entropy of a simple system
*(7>) = f
/0
(13.14)
r 1 , JT
*(T,p)
These integrals must be convergent, a requirement which could
not be satisfied, if c p and c v had a finite value for T = (because this
would make the integrals logarithmically infinite). We must con
clude, therefore, that the limiting values of c p , c v vanish
lim r o c p = Hm r . c* * 0. (13.15)
The specific heat vanishes in the vicinity of T 0: It should be
noted that Nernst's postulate is not needed for this conclusion in its
230 TEXTBOOK OF THERMODYNAMICS XIII 92
entirety: it is derived merely from the assumption that the entropy
is not infinite in the limit T ~ (lim r . S ^ )
As far as the second law is concerned, the terms /i(fl),/2(/>) in
(13.14) may be any functions of volume and pressure (section 26).
However, if we apply to them the conditions (13.08) and (13.13)
flowing from Nernst's postulate we find dfi/dv = 3/2/3/> = 0. These
terms turn out to be independent of v and p and are, therefore, constant
T fT
%dT + s , *(P,T) = *
* '0 *
dT + s , *(P,T) = *dr + s . (13.16)
SQ, the limiting value of the entropy at the absolute zero point, is a
constant independent of the process by which this point is reached, as
may be inferred from the requirement (13.07). From these expressions
can be obtained the molal thermodynamic potential, for a condensed
simple system,
PVT f
*/n
c p dlog r TSQ. (13.17)
Exercise 101. Prove eq. (13.15) in a different way. The condition lim^oS 5^ <*>
can also be written in the forms limro (W/dT)v ^ and limro (d&/dT) p
Derive from the eqs. (5.56) and (5.54) by the method used in obtaining (13.06)
r n
)  limr_o I 
or
Hmro Cv = Hmro C p = 0.
92. Planck's formulation of Nernst's postulate. It is obvious
that the last result can be generalized for the case of systems which
are not simple but depend on any number of variables. Let us sup
pose, in fact, that the zero point of temperature is attained at definite
values of all the other variables and that the entropy is found to be
equal to 5 . According to the postulate (13.07) or to eqs. (13.11),
further changes in the variables (leaving the temperature constant)
will fail to have any effect upon SQ. The state T = is, therefore, a
state of constant entropy.
Planck goes beyond Nernst in that he makes an additional assump
tion about this limiting value of the entropy: he postulates that the
entropy in the state T = is equal to zero, in every system occurring in
nature:
lim r . 5 = 0. (13.19)
XIII 93 NERNSTS HEAT POSTULATE 231
In the special case of simple systems, the entropy expressions
(13.14) are reduced to
/T /*T
^dT, s(T,p) = / C *dT. (13.20)
1 A) *
The modern statistical theory has borne out Planck's contention.
When we refer to Nernst's postulate in the following chapters, we
shall always mean Planck's formulation of it, unless expressly stated
otherwise.
Equation (13.19) helps to determine the constants or arbitrary
functions of integration in many thermodynamical relations. Thus,
the presence of the functions /(/>), f\(v) in eqs. (5.55) and (5.57) is but
a consequence of the indeterminacy of the entropy inherent in the
second law. This is brought out, for instance, by partial integration
(taking into account the formulas 3.12 and 3.26)
The parentheses are equivalent to the entropy expressions (13.14).
Replacing them by (13.20) we obtain
* = T / U v d(l/T), *  T I X p d(l/T). (13.21)
JQ ^o
Since (3^/3 F)r = p, the first of these equations gives
Knowing the internal energy as a function of T and 7, down to
T = 0, we can deduce the equation of state.
Exercise 102. The molal heats c p (in cal mol~ l , at 1 atm) of diamond, graphite,
and copper are given in the following table:
T ....... 20 K 30 40 50 60 70 80 90 100 150 200 250 300
Diamond ............................ 0.01 0.06 0.33 0.60 1.00 1.50
Graphite 0.02 0.06 0.10 0.14 0.18 0.24 0.30 0.33 0.38 0.77 1.2 1.4 2.05
Copper.. 0.1 0.4 0.8 1.4 1.94 2.5 3.0 3.5 3.9 5.0 5.4 5.7 5.9
Calculate in a rough, graphical way the absolute entropies of these substances
at 300 K.
93. Einstein's formula for the specific heat of solid bodies. In
order to establish his postulate it became necessary for Nernst to test
whether its consequences are borne out by experiment. In particular
232 TEXTBOOK OF THERMODYNAMICS XIII 93
he carried out with his pupils l an extensive program of measurements
on specific heats of solid bodies at low temperatures. In all cases the
law expressed in formula (13.15) was confirmed: as the substances
were cooled, their specific heat decreased, and, at the lowest point
attained (temperature of liquid air, later liquid hydrogen and helium),
it amounted to only a small fraction of its value at room temperature.
The previous knowledge of this question had been as follows. In
1819 Dulong and Petit deduced from their measurements the law that
elementary substances in the solid state have all the same molal (or
atomic) heat of approximately c = 6 cal/deg (section 16). How
ever, it was realized for a long time prior to Nernst's work that the
law of Dulong and Petit is not accurate and that c varies with the
temperature. A wellknown example of this was the diamond, which
exhibited, in a striking way, the type of behavior later found to be
general by Nernst and his pupils. The specific heat of diamond has
a value approaching the 6 cal/deg of Dulong and Petit only at
temperatures which are comparatively high. It decreases rapidly upon
cooling and is already pretty small (1.5 cal/deg) at room temperature,
a fact which was the source of much embarrassment for the kinetic
theory of matter.
Boltzmann had given the following simple approach to the problems
of the internal energy of elementary solid bodies: In such a body every
atom has a position of rest about which it carries out thermokinetic
oscillations. As a first approximation, one can regard the forces with
which the atom is attached to its normal position as proportional to
the distance from it. It is well known that forces of this kind (called
" elastic forces ") cause the atoms to oscillate harmonically, with a
constant frequency v. Since all atoms of an elementary body are
equivalent, each of them is attached by the same force and has the
same frequency v. What is the energy of such a model of the solid
body, according to statistical mechanics? According to (4.71), the
classical statistics of BoltzmannGibbs leads to the socalled principle
of equipartition of energy. The mean kinetic energy of every particle
subject to heat motion is kT/2 per each degree of freedom (where k
denotes Boltzmann's constant of section 4). It was also stated in
section 30 that, in the case of elastic forces, the mean potential energy
1 Monographs on Nernst 's postulate: W. Nernst, Thermodynamics and
Chemistry. New York 1907; W. Nernst, Die theoretischen und praktischen
Grundlagen des neuen Warmesatzes. Halle 1918; F. Politzer, Berechnung der
chemise hen Gleichgewichte. Stuttgart 1912. A very useful critical study of the
subject is due to W. H. Van de Sande Bakhuyzen, Het WarrateTheorema van
Nernst, Thesis. Leiden 1921.
XIII 93 NERNST'S HEAT POSTULATE 233
is equal to the mean kinetic and the total energy becomes kT per
degree of freedom. An atom is supposed to have only three degrees
of freedom and, therefore, the mean energy 3 kT. Correspondingly,
the internal energy of one mol is
u = 3n A kT + UQ = 3RT + u Qt
since WA* = R, according to eq. (1.14), and the specific heat
c = du/dT = 3R. (13.23)
Substituting R = 1.986 cal/deg, we obtain from this theory the
numerical value c = 5.95, in good agreement with the law of Dulong
and Petit. However, it does not offer any explanation for the devi
ations at low temperatures.
This difficulty was resolved by Einstein l in one of his most impor
tant papers. He pointed out that the quantum theory provides a
modification of statistics sufficient to account for these phenomena.
In fact, Planck had found that, in the case of the harmonic oscillations
of an electron of frequency v, the mean energy is not the equipartition
value kT but kTx/(e x 1), where x is an abbreviation for x = hv/kT,
and h denotes Planck's quantum of action (h = 6.55 X 10~ 27 erg sec).
Einstein made the assumption that what is true for an oscillating
electron must also apply to an oscillating atom. It was the first appli
cation of the quantum theory to atoms and molecules, and led directly
to the result
(13.24)
The entropy follows from the expression (13.20) of Nernst's theorem
in Planck s formulation
<>]
Y  log (1  e*)J. (13.26)
When the temperature T is high compared with 9 = hv/k (i.e.
x <3C 1), we find, neglecting terms of the order x
u = 3RT + UQ, c  3R, 1
(13.27)
5 = 35(1 logic), .j
1 A. Einstein, Ann. Physik 22, pp. 180, 800, 1907.
234 TEXTBOOK OF THERMODYNAMICS XIII 93
The limiting values, for high temperatures, are the same as in
classical statistics and in agreement with the law of Dulong and Petit.
On the other hand, in the vicinity of the absolute zero point (r 0,
#*oo) the formulas (13.24), (13.25) reduce to u = UQ, c = 0,
satisfying the requirements of Nernst's postulate.
Einstein's formula agrees with the general trend shown by the
experimental values of the specific heats but the numerical agreement
is not very close. Later work on the subject, which resulted in more
accurate formulas, will be treated in section 123. For theoretical uses,
however (when a fictitious solid body is as good as real one), Einstein's
model is very convenient because of its simplicity.
Exercise 103. Find the work function (^ = Ts) of the Einstein model
and calculate its approximate expressions for very high and for very low tempera
tures (x 1 and x 1).
CHAPTER XIV
BEARING OF NERNST'S POSTULATE ON THE ABSOLUTE
ENTROPY OF GASES
94. Chemical equilibrium and Nernst's postulate. The bearing of
the first and second laws of thermodynamics upon the problem of the
chemical equilibrium of perfect gases was discussed in Chapter VIII.
It was stated in section 51 that the first step in solving this problem
was the mass law of Guldberg and Waage (1867)
> log p h = log K p , (14.01)
which gives a relation between the partial pressures pi, . . . p<r of the
gases taking part in a reaction
^jV h G h = 0, (14.02)
h
conducted at a given temperature.
The next big advance was due to Gibbs (1871) who determined
the dependence (8.16) of the mass constant K p on temperature.
Within the range where the molal heats can be regarded as constant
log K p =^ log T  % + log I. (14.03)
Provided the heat of the reaction Q p and the change of the heat
capacity AC in it are known, the constant of integration / can be
determined from the measurement of K p at one single temperature.
The formula (14.03) serves, then, to calculate the mass constant for
any other temperature. We have seen that the constant / depends
upon the entropy constant of the gases in a way expressed by eq. (8.14)
(14.04)
This shows that a theoretical calculation of the constant / is beyond
the scope of the first and second law of thermodynamics, which provide
235
236 TEXTBOOK OF THERMODYNAMICS XIV 94
no way of determining the absolute value of the entropy. The formula
(14.03) with an empirical constant / represents, therefore, all the
information which can be obtained from the older thermodynamics,
with respect to the chemical equilibrium of gases. It was pointed
out in section 59 how Van t'Hoff (1878) succeeded in generalizing this
formula (still with the empirical constant /) for the case of the chem
ical equilibrium of solutes in a dilute solution.
The problem remained in this state until 1907 when Nernst brought
his postulate to bear on it. It is possible to establish a relation
between the entropy constants of a substance (ft) in its gaseous state
soh and its condensed state W c) . In fact, let us consider the equilibrium
between the vapor and the solid phase. According to eq. (7.02), the
condition of equilibrium is * = gh<Ph (c \ where <p h without upper index
represents the molal thermodynamic potential of the gas and ^ (c) that
of the condensed phase while g h = M*/W C) = v h (c) /v h is the ratio of the
molecular weights in the two phases. We suppose that the temper
ature and pressure are such that the vapor can be considered as a
perfect gas. In this case we can substitute for ^ the expression
(8.08), (8.10) where we consider c ph as constant, and for <(>h c) the eq.
(13.17) and obtain a relation which may be written in the form
<*r + iog4, (14.05)
where
), (14.06)
denotes the latent heat of vaporization per 1 mol of the vapor, and
(14.07)
is a constant called by Nernst the chemical constant of the gas.
Equation (14.05) is, of course, only the integrated form of eq. (7.05)
of section 44, It is, however, important for us to know that the
constant of integration ** is connected with the entropy constant by
the relation (14.07) which permits us to express $o* in the form
SOH = Ri h + ft** 1 *. (14.08)
Substituting this into (14.04) we find
R log I = #2>A*A Z'A^W* (14.09)
The second sum on the right side is of a structure familiar to us:
It represents the limiting value (for T = 0) of the entropy change
A5 in the reaction SJ>A (C) (?A (C> =0; in other words, it is the entropy
XIV 95 ABSOLUTE ENTROPY OF GASES 237
change which would result if all the components taking part in the
gas reaction (14.02) were condensed and the reaction conducted in the
condensed state in the vicinity of T = 0. We know from Nernst's
postulate in the form (13.07) that
Z)* W W* = lim ro AS = 0.
A
Therefore the relation (14.09) is reduced to
log /*]*. (1410)
A
This formula represents the third important advance in the
problem of the chemical equilibrium of gases and is due to Nernst.
While before 1907 the constant / had to be determined experimentally
for every separate gas reaction, eq. (14.10) made it possible to compute
it from the chemical constants i h of the components if they were known.
It was sufficient to determine by measurement the constants 4 of the
individual gases by investigating their equilibrium with the condensed
phases in order to know the constants / for all possible reactions
between them.
We have derived the relation (14.10) using the more restricted form
(13.07) of Nernst's postulate. If we accept Planck's formulation of
it (13.19), we have to put W c) = 0, and this reduces eq. (14.07) to
4 = SM/R. (14.11)
From this point of view the chemical constant of a perfect gas is,
simply, proportional to its entropy constant. Nernst's relation
(14.10) is then an immediate consequence of the definition of the
constant / given by eq. (14.04).
95. The SackurTetrode formula for the chemical constant. The
experimental determination of the chemical constant i involves
very difficult measurements. It was, therefore, fortunate that in
1912 Sackur l and Tetrode 2 succeeded in making an important step
beyond Nernst. Working independently they gave theoretical expres
sions for the chemical constant of monatomic gases which differed but
slightly. Tetrode's formula proved to be the correct one
1 0. Sackur, NernstFestschrift, p. 405, 1912; Ann. Physik 40, p. 67, 1913.
1 H. Tetrode, Ann. Phyaik 38, p. 434; 39, p. 255, 1912.
238 TEXTBOOK OF THERMODYNAMICS XIV 95
Of course, this result could not have been obtained by thermo
dynamical reasoning: the resources of thermodynamics, including
Nernst's theorem, are exhausted with the formula (14.10). Sackur
and Tetrode used statistical methods about which we shall say more
in the next section.
The argument of the logarithm in eq. (14.12) is not a pure number,
so that the numerical value of i depends on the system of units used
for the description of the gas. The reason of this lies in the structure
of eq. (14.05), where in addition to two terms of dimension zero, we
have two logarithmic terms. Taking into consideration that, for a
monatomic gas c p /R = f, the combination of the two logarithmic
terms is log (p/T^). This term depends on the units in which the
pressure and the temperature of the gas are measured and, therefore,
the same is true with respect to the chemical constant i. If we
substitute into (14.12) the numerical values in the G.C.S. sys
tem R = 8.315 X 10 7 erg/deg, n A = 6.061 X 10 23 mol" 1 , h =
6.554 X 10~ 27 erg sec, we obtain
j = i  f = 10.171 + f log M , (14.13)
an expression which corresponds to pressures measured in dyne/cm.
For use with decimal logarithms, this expression must be divided by
log 10 = 1/0.4343
jio  no ~ 1.086  4.417 + f logio /* (14.14)
If we wish to measure the pressure in atmospheres instead of
dyne/cm and to use decimal logarithms, we have to subtract
logio 1 013 249 = 6.0057, obtaining the constant
jio, atm = iio, *tm ~ 1086 =  1.589 + f lo glo M (14.15)
It was later found that the SackurTetrode expression gives the
main part of the chemical constant but is incomplete: there are
additional terms depending on the electronic and nuclear spins. The
agreement of the theoretical values with the experimental is, on the
whole, very good but it will be better to postpone the comparison
until section 120 where we shall derive the missing terms.
When accuracy is not required and the question is only as to the
order of magnitude, it is permissible to neglect the refinements and to
use the above expressions without correction. Therefore, it will be
useful to write out here the explicit form which the equation (8.16),
XIV 95 ABSOLUTE ENTROPY OF GASES 239
(8.17) of equilibrium of perfect gases assumes when the numerical
values of the constants are substituted in the form (14.15)
logio K p = ]C fc logio ph = 0.503 v f 2J vh logi
lo glo
The pressure p is here expressed in atmospheres, the heat of reac
tion Q p in cal/degree.
96. Theoretical derivation of the SackurTetrode formula.
Although Sackur and Tetrode obtained the correct expression (14.12)
of the chemical constant, the method of reasoning they used did not
stand the test of later criticism, and cannot be considered as altogether
valid. However, a rigorous derivation of this expression was given by
O. Stern 1 in 1913. As this problem lies beyond the pale of pure
thermodynamics, he had to appeal to the principles of statistics and
of the quantum theory. But the statistical elements in Stern's
argumentation are so simple that it can be given in its completeness
even in a textbook of thermodynamics. The fundamental idea is to
consider a vapor obeying the law of monatomic perfect gases in
equilibrium with its condensed phase, and to calculate its pressures,
on one hand, thermodynamically from eq. (14.05), on the other hand,
statistically. The comparison of the two results leads to the expression
for the chemical constant. This constant determines the entropy of
the perfect gas and, therefore, it can depend only on the properties
of the gaseous phase of our system remaining the same no matter
what kind of condensed phase is in equilibrium with it. It is, there
fore, permissible to consider the equilibrium even with an imaginary
condensed phase which does not really exist but could exist, inas
much as it is constructed in agreement with the laws of nature. Stern
takes as the condensed phase Einstein's model of the solid body
described in section 93 and assumes that the temperature T of the
system is sufficiently high compared with = hv/k to use the approx
imations (13.27). 2 We substitute them into eq. (14.05), taking the solid
body also as monatomic (g = 1) and its volume as negligible (v 0)
and recalling that for a monatomic perfect gas u 3.RZV2 + wo,
c p = 5.R/2. If we drop the subscript
log = f + f log T  3[1  log (hp/kT)]  "~^ + *
.    log T + 3 log (hv/k)  U ^ (C) + i. (1417)
* O. Stern, Phys. Zs. 14, p. 629, 1913; Zs. Electrochemie 25, p. 99, 1919.
pifferently from the preceding section, v is here the freqnency.
240 TEXTBOOK OF THERMODYNAMICS XIV 96
With this result the thermodynamic part of the investigation is
concluded and we proceed to the statistical part. We have specified
that the temperature is high, and this permits us to use the principles
of statistics in their traditional form since, in this case, the differences
between classical and quantum statistics disappear. All we need of
the quantum theory is already contained in eq. (14.17), which embodies
the properties of a quantized solid. In Einstein's model every atom
of the solid body is regarded as a harmonic oscillator attached to the
position of rest by a force F proportional to the distance r from it.
The equation of motion of an atom is, then,
where M is its mass and co = 2irv its angular frequency. The right side
of the equation represents the force F, so that the potential energy of
an atom at the distance r from its position of equilibrium is
e r =  f F dr = Mu 2 r 2 /2. (14.18)
/Q
These assumptions were sufficient for the theory of specific heats
(section 93) but, in order to calculate the equilibrium of the solid with
its vapor, we must specify them a little farther. How far does the
field of the force F extend? The atoms are all considered as indepen
dent in their oscillations so that it cannot reach as far as the next atom.
To satisfy this requirement we imagine, around every position of rest
as a center, a sphere of the radius a. Inside each sphere there exists a
radial force F = u?Mr while the space outside the spheres is field
free. The model of the solid body is reduced to a number of little
spheres with fields of forces in them. The atoms which are inside the
spheres belong to the condensed phase, while the free atoms outside
form the vapor.
Let dr be an element of space in one of the spheres situated at the
distance r from the center, and let Zrdr be the mean number of atoms
in this element (i.e. the mean of many observations taken at different
times). The only proposition from statistical mechanics which we
need is Boltzmann's principle (4.69) which tells us that the mean
numerical density ZT is proportional to an exponential function of the
potential energy e.
(14.19)
C being a constant of proportionality.
XIV 96 ABSOLUTE ENTROPY OF GASES 241
As the space outside the spheres is fieldfree, the potential energy
eo of the free atoms is the same as that of an atom at the border of the
sphere eo = Mw 2 a?/2. Denoting the numerical density (i.e. number
of atoms per unit volume) of the free atoms by ZQ we can apply
Boltzmann's principle also to them
0o = Cexp (eo/^r), (14.20)
whence eliminating the constant C from the two equations
z r = 2 exp
If we integrate z r over the whole volume of the sphere, we obtain
the mean number of atoms in it. It is obvious that this number must
be equal to 1. In fact, in a real solid the number of atoms is equal to
the number of positions of rest. If a part of the solid is vaporized the
positions of rest disappear with the atoms. We can write, therefore,
/FO /
Zrdr = 4irzo0f I
(14.21)
In the thermodynamic eq. (14.17) small quantities of the order
x = hv/kT were neglected. To be consistent we must neglect terms
of the same order also in the expressions derived from statistics.
This remark shows the way of getting rid of the radius a. Since the
Einstein model is not a real but an imaginary solid, the quantities v
and a can be selected at our discretion and we can choose them so that
y = Mu?a?/2kT will be still large at the high temperature at which
x = hv/kT is very small, so that exp( y) can be made smaller than
x. The integral from to a in (14.21) can then be replaced by the
integral from to oo , since the two differ only by terms of that order.
We obtain thus
1 = so(27r*r/Mco 2 )* exp (eo/r).
The potential energy of 1 mol of the gaseous phase is n A eo. It
is the energy which would have to be expended to make n A atoms free,
if all the bound atoms were at rest in the centers of their respective
spheres. This state of affairs prevails at the absolute zero of temper
ature when the atoms have no kinetic energy. The physical meaning
of n^eo is, therefore, the difference in the zero point energies of the
gaseous and the solid phases: w^eo = wo W c) . Therefore,
exp [(W c)  u )/RT] t (14.22)
242 TEXTBOOK OF THERMODYNAMICS XIV 96
since 1? = kn A and o> = 2irv. The number ZQ defines the vapor pres
sure according to eq. (1.15), p = zokT. Substituting ZQ from (14.22)
and taking the logarithm,
log p = f log 2wM +  log k  3 log h 2   log T
~ (C) . (14.23)
Comparing this expression with (14.17), we find that in forming
the difference the quantities p, T, v, UQ o (c) are eliminated and we
obtain for the chemical constant i precisely the expression (14.12) of
SackurTetrode.
In the preceding discussion we have not made use of the electric
neutrality of atoms. All our considerations and formulas would apply,
therefore, just as well to electrified particles attached to their respec
tive spheres by electric forces, provided a gas consisting of such parti
cles is subject to the law of perfect gases. The slight corrections
which are necessitated by the existence of electronic and nuclear spins
as well as the theory of chemical constants of diatomic and polyatomic
gases will be dealt with in section 120.
CHAPTER XV
CRITICAL ANALYSIS OF NERNST'S POSTULATE 1
97. Unattainability of the absolute zero. In the years following
the discovery of Nernst's theorem the question was much discussed
whether it really represented an independent and new principle of
science or was contained, in some way, in the first and second laws of
thermodynamics. The relation of the isothermal !T = to the system
of adiabatics, S = const, played an important role in these discussions.
The second law not only states that the entropy is a function of the
state but implies also that it is a unique function of the state. No
system can have two different values of the entropy at the same time.
In fact, if this were possible in some state of the system, this state
would be common to two adiabatics (say, S = S\ and S = 2, where
Si > S%) representing their intersection. This leads to a contradiction
with the postulate contained in the second law that in an adiabatic
process the entropy cannot decrease: starting from the entropy Si
one could lead the process to the state of intersection and thence along
the adiabatic 52. It follows that the system of adiabatics S = const
is a family of surfaces which do not possess intersections or envelopes.
However, it must be remembered that the adiabatic process is defined
as one in which no heat is imparted to the system, its equation being
dQ = TdS  0. (15.01)
In addition to the solution dS = 0, or S = const, this equation
has also the singular solution T = which represents at the same
time an isothermal and an adiabatic process. Taking the (/>, V)
diagram of a simple system as an example, one is inclined, at first
sight, to expect conditions as in Fig. 45: the lines of constant entropy
form a nonintersecting family (of which we give only the two members
5 = Si and 5 = 2), while the curve T = cuts across. It seems
possible to lead the system by a succession of reversible adiabatic
1 With the kind permission of Yale University Press parts of this Chapter were
patterned after the exposition by P. S. Epstein, Commentary on the Scientific Writ*
ings of J. W. Gibbs, Article 0, Sections 6, 7.
243
244 TEXTBOOK OF THERMODYNAMICS XV 97
processes from the state A to the state D of a different entropy, and
this would be a contradiction with the second law of thermodynamics.
On the other hand, this contradiction would not exist if the curve
T = were identical with one of the lines S = const, instead of being
a stranger in their midst and cutting across their system.
There are, however, two reasons why it is not permissible to
conclude from this argument that Nernst's postulate is a consequence
of the second law. The first of them was realized early, being brought
home forcibly by the example of the perfect gas which obeys the
second law but contradicts Nernst's theo
rem. In the (p, v) diagram of the perfect
gas the line T = is represented by the
axes p = and v = which serve as
asymptotes both of the system of isother
mals pv = const and of the system of
adiabatics pv y = const. For this reason
the line T = cannot be reached by any
FIG. 45. State T  in finite and reversible adiabatic process so
relation to adiabatics. that the conditions illustrated by Fig. 45
do not apply. It is true that, in the case of
the perfect gas, the isothermal T = also belongs, mathematically
speaking, to the system of lines of constant entropy, S = const, being
its extreme member S = <*> in view of the entropy expression (4.18)
5 = c p log T  R log p + *o, (15.02)
(where we suppose p j& 0, v = 0). Nevertheless, this system does not
satisfy Nernst's theorem which requires that no change of entropy be
possible for T = 0. By changing the pressure from pi to p2 we can
change s in formula (15.02) by the finite amount As r0 = R log (pi/p2),
while it retains its singular value oo .
The example of the perfect gas is a conclusive proof that the second
law of thermodynamics does not contain Nernst's theorem in its en
tirety. But attempts continued to show, at least, that certain conse
quences deduced from it could be also obtained from the second law.
In this connection there was advanced by Nernst the principle of the
unattainability of the absolute zero of temperature. 1 If one of the states
T = could be reached by a reversible adiabatic process, the diagram
of the Fig. 45 would be an adequate illustration of the actual conditions.
Pointing out that the succession of adiabatics AB, BC, CD could be
1 W. Nernst, Die theoretischen und praktischen Grundlagen des neuen WSrme
eatzes. Halle 1918.
XV 98 CRITICAL ANALYSIS OF NERNSTS POSTULATE 245
used to decrease the entropy of the system, Nernst concludes that
the assumption of attainability is in contradiction with the second law.
This conclusion was challenged by Einstein, 1 who questions that the
part BC of the process (corresponding to T = 0) can be carried out.
A reversible process is only an ideal, and there is in reality always a
certain degree of frictional waste of work, developing heat. However,
the slightest amount of heat production would throw the system off the
curve T = 0. To bring out the gist of Einstein's objection, we may
amplify it as follows. From the point of view of physics, a process can
only then be considered as defined and meaningful when an experimen
tal procedure is given by which it can be carried out. In general, the
adiabatics and isothermals satisfy this requirement. In the case of a
simple system, for instance, the adiabatic process is carried out by
compressing (or expanding) the system while it is enclosed in a heat
insulating envelope, the isothermal process by the same procedure
while the system is in thermal contact with a heat reservoir. The
particular process T = has the peculiarity that, while it is isothermal,
no heat is transferred in it to the reservoir. Nernst treats it, therefore,
as an adiabatic process. This is equivalent to dropping the heat
reservoir altogether and imagining the system adiabatically insulated
in all three branches of the Fig. 45. However, the lack of experimental
differentiation between the compressions along the lines BA and BC
makes the branch BC meaningless. In fact, no experimental direction
is given (or, indeed, possible) to insure that a compression, starting
from the point B, should cause the system to move in the curve BC
and not in BA. This is the second reason why the argument making
use of the diagram of Fig. 45 is fallacious.
These objections refer, however, only to the attempt to derive the
principle of unattainability from the second law. If Nernst's postulate
is accepted, unattainability becomes a matter of course since it states
(in Planck's formulation) that T = coincides with 5 = 0. The
process T = belongs, therefore, to the family S = const and is
identical with its extreme member. As the adiabatics S = const do
not intersect, it is obvious that 5 = cannot be reached by any
reversible process represented by one of the other adiabatics.
98. Solutions, supercooled liquids, and the statistical interpreta
tion of Nernst's postulate. Of great practical and theoretical import
tance is the question whether Nernst's postulate is a general law or
admits exceptions for certain condensed systems. Some controversy
1 A. Einstein, Structure de la matiere (Second Solvay Congress of 1913). Pub
lished 1921.
246 TEXTBOOK OF THERMODYNAMICS XV 98
arose with respect to solutions : Planck 1 maintains that his formula
tion of the theorem applies only to chemically pure substances while
in the case of a solution or mixture one should add to the entropy the
expression
so =  R(Ni + . . . + #/)/, log *,, (15.03)
*i
where xi, #2, represent the mol fractions of the components of the
system and N\, N2> their mol numbers. This term represents the
entropy of mixing the components in the case of a mixture of perfect
gases (compare section 50). According to Planck, a condensed solu
tion or mixture should have the same value of the entropy constant
since it can be converted into the gaseous phase by vaporization.
In this country Planck's views were advocated by Lewis and
Gibson. 2 These authors went beyond Planck and questioned, on
theoretical and experimental grounds, the applicability of Nernst's
theorem to supercooled liquids even of a chemically homogeneous
constitution. On the other hand, Nernst 3 had always claimed the
general validity of his theorem and was supported by other authors. 4
The contradiction was resolved, and partly reconciled, by an investiga
tion of O. Stern's 6 concerning the entropy of mixed crystals.
It will be best to say here a few words about the statistical aspect
of the problem, as both the nature of the difficulty and its ultimate
resolution are reduced in it to its simplest terms. We have seen (in
section 30) that in statistical mechanics the entropy of a state is inter
preted as proportional to the logarithm of its probability 5 = k log P f
which, in turn, is measured by the number of ways in which the state
can be realized. If we consider a chemically homogeneous crystal at
T = 0, every one of its Z atoms has a perfectly definite position of rest.
We have discussed this case in section 30 and seen that two probability
definitions are possible. If we regard permutations of identical atoms
as independent realizations of the crystal (specific definition), we obtain
P. = Z! and So = k log ZL On the other hand, if (denying the possi
bility of telling whether the atoms were permuted or not) we consider
all the permutations as one single realization (generic definition), we
find P = 1 and So = 0. Nernsfs theorem has, therefore, the generic
probability definition as its statistical equivalent.
1 M. Planck, Thermodynamik (6th ed), p. 285, 1921.
* Lewis and Gibson, J. Am. Chem. Soc. 42, p. 1542, 1920.
1 W. Nernst, Sitzungsberichte Berlin, p. 972, 1913.
E. g. W. H. Keesom, Phys. Zs. 14, p. 665, 1913.
1 0. Stern, Ann. Phyaik, 49, p. 813, 1916.
XV 98 CRITICAL ANALYSIS OF NERNST'S POSTULATE 247
For a mixed crystal consisting of atoms of two kinds (Zi and Z* in
number) the generic definition gives the probability by formula (4.66).
By the use of the Stirling approximation (log Z! = Z log Z) and of
mol fractions, xi = Z\/(Z\ + Z*), etc., it can be written as
Pi 2  xr Zl x 2 ~ z *, (15.04)
so that the zero entropy of the mixed crystal appears to be of Planck's
form
5 W ~ R (Ni + N 2 )(xi log xi + x 2 log * 2 ), (15.05)
since kZi = RNi, according to (1.14). If we call each of the realiza
tions of the mixed crystal (differing by the arrangement of at least
one pair of atoms of the two kinds) a modification of it, the number of
modifications is identical with Pi2 and we meet, apparently, with the
same difficulty which was encountered by Planck. However, the
difficulty is removed by the following remark. The number of possible
modifications only then measures the probability in a meaningful way
if all these modifications belong to the same thermodynamical state.
It was pointed out by Stern that this is not always the case. The
rearrangement of atoms changes in a slight degree the internal energy
and the other thermodynamical functions of the mixed crystal. Every
modification has, therefore, a slightly different thermodynamical
potential.
Suppose now that we have a very large number Z of copies of the
same mixed crystal at the temperature T. Owing to the thermokinetic
motions, the atoms will be continually rearranging themselves, chang
ing the modifications to which the crystals belong. We may ask,
therefore, how many of these copies are, at any given time, in the
modification j having the thermodynamic potential $/, and we shall
show in the next section that their number is
Z y = constexp (*,/W). (15.06)
Correspondingly the ratio Z//Z represents the probability of find
ing a given crystal in the modification j. Since the differences in the
function $, are very minute, all the numbers Z/ are practically equal at
all temperatures which are not too low. This means that all the
modifications are equally probable representing statistically equiva
lent realizations of our crystal. It is appropriate to count, as the
probability of the crystal, the number of these realizations, as was
done in formula (15.04), and this leads to the expression (15.05) for
the entropy of mixing. On the other hand, at very low temperatures
even very small differences in */ begin to tell because of the factor.
248 TEXTBOOK OF THERMODYNAMICS XV 98
1/r in the exponent. The nurfibers Z, become unequal and, in the
extreme case of the immediate vicinity of T = 0, the modification j m
with the lowest thermodynamical potential $ m i n dominates to such an
extent that all the others are impossible. In the vicinity of the abso
lute zero the mixed crystal is, therefore, in the modification j m . Since
every modification can be realized in only one way, the probability
of the crystal, instead of (15.04), becomes Pi 2 = 1 and the zero
entropy So = 1, in agreement with Nernst's postulate. Offhand it is
conceivable that several modifications could have the same thermo
dynamic potential, and its lowest value $ mln might be shared by n of
them. In this case, the probability at T = would be Pi 2 = n and
the entropy So = k log n. 'The purport of Nernst's postulate is, there
fore (as far as mixed crystals are concerned), that there is one and only
one modification of lowest thermodynamic potential.
Similar considerations apply to the entropies of supercooled
liquids. In a liquid the positions of the atoms and molecules are
irregular and the possibilities of arranging them more numerous than
in a crystal. This makes their entropy, at high temperatures, larger
than that of a crystal. At low temperatures, however, the regular
crystalline modification is the most probable and, in the vicinity of
r = 0, even the only possible arrangement of atoms.
From the theoretical point of view Nernst's postulate is, therefore,
completely vindicated since there is no evidence that it admits of any
exceptions. However, the differences in $/ on which the above argu
ment depends are so very small that one can expect a perfectly regular
arrangement only at extremely low temperatures which are, in many
cases, below the range accessible to our experimental technique.
Another important consideration is the time element since, under these
conditions, all processes are extremely slow. For this reason one can
not be certain that the substances under investigation have reached
the state of true thermodynamic equilibrium. In this sense, we may
say that the view of Planck and of Lewis and Gibson is justifiable from
the practical standpoint of experimental measurements. This is, in
fact, borne out by the recent and very accurate observations l of
Giauque and his collaborators. The method used by these authors
is based on the fact that the absolute entropies of many diatomic and
polyatomic gases can be calculated with great accuracy from spec
troscopic data (compare section 119). They take the same sub
stances in the solid state at very low temperatures and measure the
1 Older work pointing in the same direction was due to Gibson, Parks, and
Latimer (J. Am. Chem. Soc. 42, p. 1542, 1920) and Gibson and Giauque (ibidem
45, p. 93, 1923).
XV 99 CRITICAL ANALYSIS OF NERNST'S POSTULATE 249
entropy difference attending their gradual transformation into gases
by heating and subsequent vaporization. In this way it was found
that hydrogen, 1 carbon monoxide, 2 nitrous oxide, 3 and ice 4 possess
still appreciable entropies as solids at about 15 K. The lattice
theory of solid hydrogen was treated by Pauling, 5 who showed
that the crystals are built up of para and orthomolecules (com
pare section 118) whose distribution is entirely irregular at tem
peratures above 5 K. The interpretation which Giauque and
coworkers give with respect to the other substances just mentioned is
that their molecules are not uniquely oriented but even at temperatures
as low as 15 K have still the choice between several orientations.
Pauling 6 investigated the case of ice and finds the entropy R log (3/2)
= 0.805 cal deg" 1 mol" 1 , due to the statistical weight of the indeter
minacy of orientation. This is in excellent agreement with the
experimental value 0.82 cal deg 1 mol"" 1 found by Giauque and Stout.
99. Equilibrium of modifications of a mixed crystal. There remains
to prove the formula (15.06) on which the conclusions of the preceding
section rest. We shall do this by an argument which is thermo
dynamical in its main points. Its essential idea (but not its mathe
matical form) is due to Stern. The formula in question refers to the
equilibrium of a very large number of copies of the same mixed
crystal. In order to deduce it, we must provide an ideal experimental
arrangement by which such an equilibrium can be secured. Let all
these crystal copies be contained in an enormously large vessel where
they are floating in a neutral gas medium removed from the action of
gravity. Through their Brownian movements they will set themselves
into thermal equilibrium with the medium and with one another.
Provided the size of the crystals is small compared with their mean
distance, such a suspension can be regarded as a perfect gas with
extremely large molecules obeying all the laws valid for gases (compare
section 146). In our particular case every modification of the crystal
may be considered as a separate gas, so that we have a mixture of as
many gases as there are modifications.
Any modification (1) can be converted through the internal
rearrangement of its atoms into any other modification (2), and we can
apply to this process the laws of chemical equilibrium of perfect gases.
1 W. F. Giauque and H. L. Johnston, J. Am. Chem. Soc. 50, p. 3221, 1928.
1 J. O. Clayton and Giauque, ibidem 54, p. 2610, 1932.
R. W. Blue and Giauque, ibidem 57, p. 991, 1935.
4 Giauque and Ashley, Phys. Rev. 43, p. 91, 1933; Giauque and Stout, J. Am.
Chem. Soc. 58, p. 1144, 1936.
L. Pauling, Phys. Rev. 36, p. 430, 1930.
L. Pauling, J. Am. Chem. Soc. 57, p. 2680, 1935.
250 TEXTBOOK OF THERMODYNAMICS XV 99
In the law of reaction (6.40) we have to substitute v\ = 1*2 = 1,
obtaining
Gi  G 2  0, (15.07)
while the equation of equilibrium (6.43) is reduced to
<f>i  <p 2 = 0. (15.08)
The difference between the system under consideration and the
ordinary perfect gases of Chapter VIII lies in the fact that in the
present case the thermodynamic potential consists of two parts
9 = n + v (15.09)
<pi is the " internal " potential of the crystals themselves, <p e the
" external " part due to the Brownian motions. For <p e we have to
use the expression (8.08) of the thermodynamic potential in perfect
gases, omitting the terms UQ TSQ which refer to the intrinsic proper
ties of the " molecules " (here crystals) and are, therefore, included in
<pi. The total potential is, therefore,
V = T(R log p  c p log T + c p ) + Vi. (15.10)
At this stage of the reasoning we have to invoke another result of
statistical mechanics, the fact that the " external " molal heat c p (due
to the movements of the crystals) depends only on the number of
degrees of freedom of the " molecules " regarded as rigid bodies. It is,
therefore, the same for all modifications: c p \ = C P 2. In substituting
(15.10) into the condition (15.08), the terms with c p will cancel out
leaving
RT log (Pi/fr) =  (<pn  ^2). (15.11)
<pn is the molal thermodynamic potential, i.e. it is referred to n A
crystals (n A being the Avogadro number). It stands, therefore, in the
following relation with the thermodynamic potential <>, of an individual
crystal used in the preceding section <pn = n A $,. Noting that R = n A k,
we obtain
log (Pi/fr) (*i *2)/*r. (15.12)
pi and p2 are the partial pressures in our vessel of the two gases con
sisting of the modifications (1) and (2). These pressures are propor
tional with the numbers Zi and 2 in which these modifications are
represented in the vessel, according to the formula (1.15). The last
equation becomes, therefore, identical with (15.06), and this is what
we set out to prove.
XV 99 CRITICAL ANALYSIS OF NERNST'S POSTULATE 251
We have obtained this result by applying to our imaginary sus
pension the laws of the classical perfect gas. It may be objected that
these laws have doubtful validity at very low temperatures. How
ever, we shall see in the next chapter that the approximation given by
them is the better, the larger the weight of the molecules. Since, in
our ideal experiments, we may attribute to the crystals any size, the
accuracy may be considered as of any desired degree even in the
immediate vicinity of the zero point.
CHAPTER XVI
DEGENERATE PERFECT GASES
100. Equation of state of the monatomic degenerate gas. Nernst's
theorem was enunciated only for condensed systems, and it was
realized, from the beginning, that the classical perfect gas fails to
fulfill it. In fact, the entropy of the perfect gas has the form (4.18)
which would need an infinite entropy constant in order to meet the
requirement s r _ The coefficient of expansion of the perfect
gas is, according to (1.19), a. = 1/T, while the specific heats c v and c p
are constants independent of temperature (section 15), in contradic
tion with the two immediate consequences (13.09) and (13.15) of the
theorem which would require lim r . a = and lim r . c = 0. Begin
ning with Sackur and Tetrode (section 95) there were many attempts
to subject the perfect gas to quantum conditions and to obtain a cor
rected equation of state. As long as these endeavors were based on
the classical statistics of BoltzmannPlanck they failed, but the
advent of new types of statistics led to a great success, the discovery
of the correct laws of perfect gases which happen to be in agreement
with Nernst's theory.
A textbook of thermodynamics is not the place to enlarge upon the
statistical principles underlying the new theory of the perfect gases.
We shall give only a brief reference to questions of statistics in section
104 of this chapter and shall take the equation of state as given, in the
same way as we did not inquire into the statistical justification of the
classical perfect gas. If we restrict ourselves to monatomic gases, the
corrected equation of state can be written in the form
(16.01)
It differs from the old equation of perfect gases (1.13) only by the
factor P(6)/9 appearing on the right side. By P(G) is meant a certain
function of the argument
,,,
252
XVI 100 DEGENERATE PERFECT GASES 253
where the letters have the same meaning as before. In particular,
/x is the atomic weight, h Planck's quantum of action (6.554 X 10 ~ 27
erg sec), and n A the Avogadro number. The function P(9) is so com
plicated that it cannot be given by an explicit formula but only
implicitly. In order to define it we introduce two auxiliary functions
F(A) and G(A) of a new argument A by the equations
r ^'
l
(16.03)
Ax"e*dx
1 dAe~*
We have to distinguish three types of perfect gases which differ,
according to the value assigned to the constant 6 and according to
the range of variability of the variable A :
(1) The classical or BoltzmannPlanck gas
6 = 0, 0^4^oo.
(2) The Fermi degenerate gas
d = 1, ^ A ^ oo.
(3) The EinsteinBose degenerate gas
+!, O^A^l.
When 5 =+ 1, A cannot be larger than 1 because otherwise the
integrals (16.03) would be divergent.
In all three cases the connection between F(A), G(A) t on one
hand, and P(0), 0, on the other, is the same, namely:
F(A) = 8*. (16.04)
This relation expresses A in terms of 9, so that G(A) can now
also be regarded as a function of 6. Therefore, the second relation
which we impose,
P(0) = 0* G(A), (16.05)
amounts to defining P(0) as a function of 0.
It should be noted that the functions F(A) and G(A) are not inde
pendent but there exists between them the relation
. (16.06)
dA A
which is readily obtained by multiplying the numerator and denom
254
TEXTBOOK OF THERMODYNAMICS
XVI 100
inator of the integral for G(A) by e*/A, then differentiating the result
with respect to A and partially integrating with respect to x.
In the range A ^ 1 the functions (16.03) can be represented by
the following series
(16.07)
The opposite case of very large values of A, (A ~2> 1) is of interest
only in connection with the Fermi gas, 5 = 1. The approximations
for the two functions are, then, as follows
57
(log ^
(16.08)
In the case (d = 1) of the EinsteinBose gas, A cannot be larger
than 1 and, for the classical gas (5 = 0), no approximation is needed
since F(A) = G(A) = A are, then, the rigorous expressions for the
functions.
It is clear from the formulas (16.07) and (16.08) that F(A) is
parallel with A in the sense of being small or large when A is small or
large. On the other hand the equation (16.04) shows that 9 is in an
inverse relation to F(A) and, therefore, to A: when A is small, is
large, and vice versa. From the four equations (16.07), (16.04), and
(16.05), we can eliminate by the method of successive approximations
the three quantities A, F(A), G(A) and obtain the following approx
imate relation between P(0) and 6:
(0 1)
The same procedure carried out with the formulas (16.08), (16.04),
and (16.05) applies only to the case d =  1 of the Fermi gas
(0 1)
XVI 101 DEGENERATE PERFECT GASES 255
We see from (16.09) that, for very large values of the variable 9,
the function P(9)/9 becomes closely equal to 1 so that eq. (16.01) takes
the classical form pv = RT. By its definition (16.02) the quantity 6
is proportional to v*T, it is large when the temperature is high or
the density low. Under these conditions the gases of EinsteinBose
and Fermi do not deviate in their properties from the classical perfect
gas: they are nondegenerate. As 9 decreases the deviation from the
classical laws becomes more and more marked : the degree of degener
ation increases. However, we shall postpone the quantitative discus
sion of the degree of degeneration until section 105.
101. The thermodynamic characteristics of the degenerate gas.
From the equation of state (16.01) the internal energy can be obtained
by means of the relation (4.23)
.
or integrating with respect to dv,
The function f(T) is determined by the condition that the gas is
nondegenerate at high temperatures. The internal energy must take,
for T >oo, the classical value (3.18) without becoming infinite for
T = 0. It is convenient to change the variable of integration from
v to 9, given by the relation (16.02) which can be also written
f log 9 = f log T + log v + const. (16.13)
Since T is to be considered as constant with respect to the inte
gration,
dv
On the other hand, T(dp/dT) v  9(3/>/99) t ,. After these substi
tutions the integration can be readily carried out, giving, with a suit
able disposal of the f unction /(r),
P(Q)
This is, in fact, identical with (3.18) in the limit 9 oo, since for
a classical monatomic gas c v = ^ R.
256 TEXTBOOK OF THERMODYNAMICS XVI 101
Knowing the internal energy u and the pressure p, we can calculate
the entropy differential (4.15), ds = (du + pdv)/T. According to the
expressions (16.01), (16.13), and (16.15) for p, v, and u, we can choose
as variables T and G. The differential dv can be expressed in terms
of dT and dQ: it follows from (16.13)
3dQ ^$dT dv
2 9 ~2 r + v 9
This change of variables reduces ds to the form
depending only on 9.
Nernst's postulate in the form (13.19) can be directly applied only
in the case of the Fermi gas. In fact, in the case of the EinsteinBose
gas the variable A cannot be larger than 1, and the relation (16.04)
shows that 9 cannot sink below a certain finite limit and is not defined
in the vicinity of T = 0. We pointed out in section 1 that the postu
late does not apply to the classical perfect gas. Therefore the validity
of the simple entropy expression
i R J. r (16  17)
is restricted to the Fermi gas. The integration can be carried
out by transforming back to the variable A with the help of
the relations (16.04) and (16.05). The first of them can be written as
log 9 = f log F(A) t or differentiating
d9 2dF
9 ~~3T'
In this way the integrand is transformed into
dP(Q) __$ G W5dG_2dG
9 ~~ 3 F 2 + 3 F 3 F*
The first two terms give $d(G/F) and we can write instead of the
last %dA/A, because of the relation (16.06). Hence
The condition 5 * 0, at the limit 9 = 0, or A = oo is taken care
of because, for very large values of A, the formulas (16.08) give
XVI 102 DEGENERATE PERFECT GASES 257
log A. We shall show by a special investigation in section
103 that the form (16.18) of the entropy expression holds also in the
case of the Einstein Bose gas.
Finally, the formulas (16.01), (16.15), and (16.18) give for the molal
thermodynamic potential <p = u Ts + pv the simple expression
<p = RT log A + u Q , (16.19)
which is true for both kinds of degenerate gases.
102. Chemical constant of degenerate gases. Their relation to
Nernst's postulate. We have seen in section 1 that for high tempera
tures or low densities (6 <3C 1 and A ^> 1) the gases of Fermi and
EinsteinBose are non degenerate, approximating in their properties
the classical perfect gas. When the variable A is extremely large, the
functions (16.07) are reduced to F(A) = G(A) = A, whence eq. (16.04)
gives A 9"^. The entropy expression (16.18), therefore, takes
the form
(16.20)
or because of (16.02) and R = n A k,
s = R log v + %R log T + R i 1
= %R log T  R log p + Ri,
(16.21)
The entropy has the classical form (4.18), specialized for the monatomic
gas (c v = !?, c p = j^R), while the constant i exactly coincides, with
the SackurTetrode expression (14.12). Putting the question as to
the validity of Nernst's postulate we must remember that we so far
obtained complete expressions for the energy and equation of state
(valid down to the vicinity of T = 0) only in the special case of the
Fermi gas. In deriving them we made explicit use of Nernst's theorem
through eq. (16.17) so that there is no question but that it is fulfilled.
It will be well, however, to check here the two immediate consequences
from the theorem which we have derived in sections 90 and 91, that
the specific heat and the coefficient of thermal expansion must vanish
for T = 0.
For the specific heat we have from (16.15) and (16.10) the value
258 TEXTBOOK OF THERMODYNAMICS XVI 103
which is proportional to T, because of formula (16.02), and vanishes
with T.
For the coefficient of thermal expansion, eqs. (1.02) and (1.05) give
Since p differs from u only by the factor 2/3v,
(16.24)
(dp\ _ 2*
\dT/ v 3 v '
so that a also vanishes for T = in a linear way.
To elucidate the conditions in the case of the EinsteinBose gas we
need a separate investigation to which we now turn.
103. Condensation of the EinsteinBose gas. It was pointed out
in section 100 that in the EinsteinBose case (5 = 1) the variable A
cannot be larger than 1. In view of this, 9 has a lower limit 9 m m
determined by eq. (16.04) with 4 = 1:
F(l) = 6. (16.25)
Now F(l) = 2.6123, and the numerical value of the limit is, there
fore, 6 m i n = 0.528. Since the parameter 9 depends on v and T 9 accord
ing to eq. (16.02), we may ask what will happen if, keeping the temper
ature of the gas constant, we should compress it sufficiently to make 9
fall below the limit 9 mln (or else decrease the temperature keeping the
volume constant). To this question Einstein gives the answer that a
part of the system will cease to be gaseous and will be precipitated out,
forming a condensed phase in equilibrium with the remaining part of
the gas. The condition of equilibrium is given by the equation (7.03)
9  W, (16.26)
where <p and <ppr represent the thermodynamic potentials of the
gaseous phase and of the precipitate. Since the expressions (16.18)
and (16.19), derived for the entropy and thermodynamic potential of
the Fermi gas, give the correct value of the chemical constant, it is
desirable to conserve them for the EinsteinBose gas. Under the
conditions of condensation (saturated state) 9 = 9 mln and A = 1, so
that eq. (16.19) gives <p = UQ. Consequently, <ppr = u pr Ts pr + pVpr
must be also equal to UQ, for all temperatures and pressures compatible
with 9 9 m in. This requires u^ = o s pr = 0, v pr = and means
that in the solid phase the material is much in the same state which
prevails at T 0. It does not possess any thermokinetic energy, and
XVI 104
DEGENERATE PERFECT GASES
259
it has a perfectly regular arrangement of atoms and a negligible
volume.
In support of Einstein's explanation we can point to the peculiar
shape of the isothermals of the EinsteinBose gas. 1 By means of (16.02)
and (16.05) we obtain from (16.01) the expression for the pressure
\ n A ) A 3
G(A),
(16.27)
and calculate from it the partial derivative
/aA _ dp t dA
\dv/T~ dA ' dF
(^).
\dV/T
Because of (16.06) and (16.04), this gives
dp\ _ RT 1 dA
at; A" v 2 Q 3 A dF
f
V
(16.28)
It is easy to see from the definition (16.03) of the function F(A)
that (in the case 6 = 1) the derivative
dF/dA becomes infinite for A = 1. Con
sequently, (dp/dv)T = 0, for A = 1, 6 = 9 m i n .
In the state of saturation the gas is, there
fore, characterized by an infinite compress
ibility and its isothermals by horizontal
tangents. These conditions are illustrated
in Fig. 46, where the solid lines represent
the isothermals and the dotted curve cor
responds to 8 = m i n . We have seen in the
theory of the Van der Waals equation that
the points (dp/d^T = are those where a
single phase of the gas begins to be completely
unstable. The condensation of the ideal gas
of EinsteinBose is, therefore, closely analo
gous to the conditions in real gases.
104. Statistics underlying the theory of degenerate gases. Rule
of alternation. 2 We have introduced and treated eq. (16.01) of
degenerate gases in a purely formal manner. It will be well to say now
a few words about its origin and about the question of which gases of
1 Compare P. S. Epstein, Commentary on the Scientific Writings of J. W. Gibbs,
Article 0, Section 10.
1 For a detailed exposition of quantum statistical principles, see: P. S. Epstein,
ibidem, Article V.
FIG. 46. Isothermals in
EinsteinBose degenerate
gas.
260 TEXTBOOK OF THERMODYNAMICS XVI 104
nature are subject to degeneration of the types of Fermi and of
EinsteinBose. Classical statistics and the statistics of the older
quantum theory were dominated by Boltzmann's principle (4.70):
If we have a single particle (or other system) which can assume dif
ferent dynamical states, the probability of finding it in a state char
acterized by the energy e, is
P i = Cexp(e,/*r), (16.29)
C being a constant. Nothing is changed if we consider Z particles of
one and the same kind under joint external conditions (field of forces
or vessel), determining a set of energy levels ei, 62, . . . e/, . . . which
any of the particles can assume. The mean number Z/ of particles
occupying the energy level e/ is again given by
Zy/Z = C exp ( e,/*r), (16.30)
because in the classical theory the particles are regarded as statistically
independent: the probability of a particle assuming any state is not
influenced by the presence of the other particles and remains the
same as if it were alone.
With the advent of wave mechanics the point of view had to be
changed. In the case of a single particle, Boltzmann's principle (16.29)
remains valid but Z particles forming a system of the kind mentioned
above can no longer be regarded as statistically independent. 1 The
primary particles of matter (the proton, neutron, and electron) fulfill
the Pauli exclusion principle : No more than one of Z identical particles
can occupy any given quantum state. (In the language of wave mechan
1 It was pointed out in section 30 that the fundamental difference between
classical and quantum statistics lies in the question whether identical particles can
be told apart. This question has an intimate relation to Heisenberg's principle of
indetermination which states that the uncertainties A* (of a coordinate of position)
and A/> x (of the conjugate momentum) are restricted by the inequality Ax &p x <h/4T.
In fact, when the uncertainty of position of each particle in a gas is small compared
with the mean distance between the particles (v/n A )^, then there exists the con
ceptual possibility of keeping track of any particle and mentally identifying it as
distinct from the others. Therefore, classical statistics must give a correct result
when the condition (/n 4 )**Ax is satisfied. On the other hand, when the two
quantities are of the same order of magnitude, there is no way of telling which
particle is which, the concept of permutation loses its sense, and classical statistics
can no longer be applied. The uncertainty of the momentum is, obviously, the
mean momentum itself A/> x J, since the particlejnay have any velocity. As long
as classical statistics (equipartition) is valid p x * M*v x * = MkT. We obtain,
therefore the condition (MkT)^(v/n A )^ A/4r, or using the notation (16.02),
6l/8r. This is the deeper reason why degeneration is absent when 6 is large
and sets in when this quantity becomes small.
XVI 104 DEGENERATE PERFECT GASES 261
ics the same thing is stated by saying that the primary particles have
antisymmetric wave functions.) We cannot enter here into the details
of the statistical calculations based on the exclusion principle and
must be satisfied with their results: they lead to the modified formula
(5 a l), which represents the fundamental law of Fermi's statistics.
The quantity A gives a measure of the deviation from the classical
conditions of the formula (16.29). It is, therefore, obvious that A must
be a function both of the particle density in the system and of its
temperature. In fact, the sum of all Zj is the total number of particles,
Z, leading to the condition
Another condition follows from the energy relations. The total
energy of the Z, particles in the state j is e,Z, and, therefore, the total
internal energy of the system U = 2e,Z, + UQ or
U  U = ZC
t
It is possible to eliminate C from these two equations and to obtain
A as a function of U, Z, and T.
Let us now turn from primary particles to composite ones, for
instance, to atomic nuclei built up of protons and neutrons. If anti
symmetric wave functions are used for the protons and neutrons
within the necleus, the translational motions of it, as a whole, are
described as follows: (a) by antisymmetric wave functions, if the
number of primary particles in the nucleus is odd, (6) by symmetric
wave functions, if this number is even. Nuclei of the first kind have
the same statistical properties as the primary particles themselves,
they obey the exclusion principle and the formula (16.31) of Fermi's
statistics (with 5 = 1). On the other hand, nuclei of the second
kind possess entirely different qualities: the statistical calculation
shows that in their case the number Z, can be expressed by the formula
(16.31) with d =+ 1 which characterizes the statistics of EinsteinBose.
The law that the statistics of Fermi applies to composite particles
with an odd number of primary elements, and EinsteinBose's to those
with an even number, is known as the rule of alternation. Of course,
this rule is theoretically valid also in the case of neutral atoms which
262 TEXTBOOK OF THERMODYNAMICS XVI 104
contain primary particles of three kinds: electrons in addition to
protons and neutrons. The experimental test of the rule could not be
carried out by observing the degeneration of gases (compare next
section) but by study of molecular band spectra whose structure also
depends on the statistical properties of the nuclei composing the
molecule. In the present state of our knowledge the rule of alternation
seems to hold without exception. Doubts exist only with respect to
the sulfur nucleus, but the experimental data are insufficient to decide
whether they are well founded. There was introduced into the theory
of nuclear structure the hypothesis of the existence of another primary
particle, the neutrino, which is supposed to have a very minute mass.
This assumption proved very useful in the theory /3ray emission and
of nuclear transformations, but it seems that the neutrino could be
dispensed with as far as the rule of alternation is concerned.
The conditions (16.32) and (16.33) determining the quantity A are
closely related to the integrals (16.03) of section 100. If the system is
a gas, contained in a vessel of the volume F, the energy levels are
determined by the vessel. It is shown in the quantum theory that in
the interval of energy values between e and e + Ae there lie
2irF(2/i/w^^ 2 )^e^Ae levels, and, if V is sufficiently large, their dis
tribution is practically continuous. By introducing the variable
x = e/kT the sums (16.32) and (16.33) can be thrown into the form
of integrals proportional to (16.03). The comparison shows
~
U  Uo = %ZkTG(A}/F(A).
(16.34)
Thus, the probability definition (16.31) leads directly to an energy
expression equivalent with the formula (16.15).
The gases here discussed are perfect in the sense that their molecules
(or atoms) have no extension and there are no forces of interaction
between them. However, the statistical limitations imposed on identi
cal particles by Pauli's exclusion principle or the requirements of
EinsteinBose have an effect upon the thermal and caloric equations
of state which resembles the effect of force interactions (compare next
section). It must be emphasized that no such limitations exist for
nonidentical particles. In a mixture of several degenerate perfect
gases, there is no mutual influence between them, and each gas behaves
as if it were alone and the others absent. This implies, in particular,
that the formula
P  Pi + p2 + ...+P., (16.35)
XVI 105 DEGENERATE PERFECT GASES 263
representing the total pressure as the sum of the partial pressures of
the component gases (each calculated as if the rest were not present)
remains valid.
105. Chances of observing the degeneration experimentally.
Some of the laws of perfect gases have the same mathematical form
for all three kinds of them (classical, Fermi, and EinsteinBose).
Such, for instance, is the equation of the adiabatic process: it follows
from the expression of the entropy differential (16.16) that s = const,
when 9 = const. According to (16.02) this can be written
const, (16.36)
or multiplying by pv from eq. (16.01)
pv* = const. (16.37)
The relation between pressure and energy density resulting from
(16.01) and (16.15)
is also valid for the classical and both degenerate gases.
On the other hand, the pressure expressed in terms of the molal
volume and temperature depends on the nature of the gas. Written
out explicitly the expressions are
p = * (i  5/2*6")
or
h?n A ** 1 \ 7
^Ukf^ V + '")'
(9 <C 1 occurs only in the Fermi gas)
'
Substituting into the expression (16.02) for 9 the numerical values
of the constants k,n A , h, and VQ (molal volume at normal conditions),
we find
9 = 3.67 M () T. (16.41)
\VQ/
This gives for helium (/*  4.00) at C and 1 atm, 9  4033, so
that the condition 9 ^> 1 is fulfilled for the permanent gases of
264 TEXTBOOK OF THERMODYNAMICS XVI 105
nature. On the other hand, electrons in metals also form a gas (com
pare next chapter) of the atomic weight /z = 1/1821 = 0.000549
and with a remarkably small molal volume, namely of the order
V/VQ 1/2000. In such an electron gas the quantity 9 has at room
temperature a value of the order of 9 = 0.004, satisfying the opposite
condition 9 <C 1. The two extreme cases represented, respectively,
by the formulas (16.39) and (16.40) happen, therefore, to be sufficient
for the treatment of all the degenerate systems occurring in nature.
For the permanent gases, eq. (16.39) must be used, and its form is
analogous to the Van der Waals equation of state (1.21) which, in the
case of not too small molal volume (v J> b), can be written
Perfect gases have, of course, no extension of molecules (6 = 0),
but the EinsteinBose gas (6=1) acts as if its molecules were attract
ing one another, the Fermi gas (5 = 1), as if they exercised repulsive
forces. This behavior is in harmony with the statistical meaning of
degeneration in the two cases (section 104).
The most promising gas for observations is helium: it is the
monatomic gas of lowest atomic weight, and, because of this, has
the highest degeneration. Moreover, it can be cooled to lower
temperatures than any other gas. But even in helium the condi
tions are not favorable, the Van der Waals terms of eq. (16.40) are
in it (under normal conditions) b/v = 1.1 X lQ*,a/RTv = 6.8 X 10~ 5
while the degeneration term of (16.39) has the much smaller value
1/2 H 9** =  6.9 X 10 ~ 7 . It is true that the ratio becomes more fav
orable at lower temperatures, but barring the immediate vicinity of
the critical state the constant 9 remains fairly large even then. The
effect is, therefore, completely obscured by the Van der Waals
forces.
The velocity of sound can be measured with considerable accuracy
even at very low temperatures and in a small volume of gas. The de
terminations of Keesom in helium at 4 K and 0.08 atm gave the value
to be expected in a nondegenerate gas. It is easy to see the reason
for this. Since the equation of the adiabatic (16.37) is the same as
for the classical perfect gas, the expression (3.40) for the velocity
of sound also remains valid. This gives (with y = ^)
 'RT (1  6/2 M 6). (16.43)
3 /* 3/*'
XVI 105 DEGENERATE PERFECT GASES 265
Under the conditions of Keesom 9 = 24 and  1 /2 W 9* =  1 .5 X H) 3 .
The effect of the degeneration is only 0.15% and entirely within the
errors of measurement.
Summarizing we can say that the observation of degeneration in
atomic gases is hopeless. The only system in which the phenomenon
is accessible to the experimental test is the electron gas which is
treated in the next chapter.
CHAPTER XVII
ELECTRON AND ION CLOUDS
106. Thermodynamic properties of charged gases. As long ago
as 1888 the pioneers of thermodynamics, Arrhenius and Ostwald, 1
applied the mass law to electrified particles (ions). We are indebted
to them and their pupils for extensive experimental data which show
conclusively that the thermodynamical theory of chemical equilibrium
applies to ions just as well as to neutral atoms. The idea that a cloud
of electrons can be also described by the laws of perfect gases proved
useful in the theory of metallic conduction and associated phenomena.
It was introduced by Drude 2 and Richardson, 3 who dealt with it from
the statistical point of view, while the first thermodynamical treatment
was due to H. A. Wilson. 4
A part of the internal energy of such a gas must be of electric
origin. Let us focus our attention on a portion of the gas so small
that we can consider the outer electric potential 12 as constant over its
volume r. The potential 2 is produced, partly, by causes extraneous
to the gas, partly by charges on the gas particles outside the volume r.
The contribution to the total electric energy of the portion within r
consists of two items: the mutual energy of this portion and of the
rest of the system, and the electrostatic energy of the volume r itself.
The first item is proportional to the first power of the volume r ; the
second goes mainly with r 2 and only rarely contains a portion linear
in r. The latter case may arise when the charges within r are of both
signs so that there is a possibility of pairing, i.e. of positive charges
hovering, by preference, in the vicinity of negative ones, and vice
versa, a phenomenon which we shall discuss in section 115. At present
we shall treat the case where pairing is either absent or negligible, so
that the inner electric energy of the volume r is quadratic in r. If we
select r sufficiently small, the quadratic term becomes negligible com
1 W. Ostwald, Zs. phys. Chemie 2, pp. 36, 270, 1888.
* P. Drude, Ann. Physik 1, p. 566; 3, p. 369, 1901.
O. W. Richardson, Proc. Camb. Phil. Soc. 11, p. 286, 1901; Philosophical
Transactions (A) 201, p. 497, 1903.
H. A. Wilson, Philosophical Transactions (A) 202, p. 258, 1903.
266
XVII 107 ELECTRON AND ION CLOUDS 267
pared with the mutual electric (mentioned above) and the caloric
terms which are both proportional to r. Let us denote the charge per
mol by / = en A (e being the charge on each particle). 1 The electric
energy of the gas in the volume r (due to the outer potential 8) will
then be/12 per mol. This is the only part of the electric energy which
has to be considered and added to the internal energy of caloric origin.
For instance, if we assume that the caloric part has the expression
(3.18), valid for a perfect gas, the internal energy becomes
u = c v T + uo+fto. (17.01)
The additional term /12 is independent of the condition of the gas
in the volume r. It is a constant with respect to it and not one of the
parameters describing its state. The parameters, subject to change,
defined within r, are temperature, pressure, and specific volume, as in
every simple system. It has been established by statistical reasoning
and by experiment that the conditions concerning the equation of
state and the caloric internal energy are in no way different in a
charged gas from those in the uncharged. The equation
pv = RT, (17.02)
together with (17.01) define the classical (or nondegenerate) perfect
charged gas and describe the phenomena in clouds of ions and electrons
to the same degree of approximation as the equations lacking the
electric term describe the ordinary permanent gases.
Since /ft must be considered as constant, with respect to changes
in the volume r, it may be regarded as forming part of the energy
constant UQ: in fact, we have seen in section 98 that UQ itself repre
sents, in part, potential energy. Therefore, the entropy differential
ds does not contain the electric potential so that we find for s precisely
the same expression (4.18) as in uncharged gases. On the other hand,
the thermodynamic potential <p becomes according to (5.41), (5.43)
<p = RT log p  CpT log T + (c p  5 )r + u + /0. (17.03)
107. Equilibrium in charged gases. The problem of equilibrium
was treated in section 40 in its generality, but the conditions in charged
systems are somewhat peculiar, and it will be necessary to say a word
how they fit into the picture there given. The phases in a system
consisting of charged gases may be determined by external conditions
and separated by surfaces of discontinuity, for instance, when we
1 f becomes identical with the Faraday F when the charge is monovalent and
positive. In general, / = <rF, if <r is the valency; for the electron (whose charge
ia negative)/  F.
268
TEXTBOOK OF THERMODYNAMICS
XVII 107
consider the free electrons inside a heated metal, as one phase, and the
electron cloud outside it, as another. However, when the electric
potential Q changes continuously with the position in the gas each
small volume r, considered in the preceding section, may be also
regarded as a separate phase of the system because the electric condi
tions in it are different from those of the rest of the gas. In either case
the potential is one of the essential characteristics of the phase, and
this fact has an important bearing on the equilibrium. In section 40,
we mentioned two possible cases, represented by eqs. (6.26) and (6.29),
arising when there is no transfer of matter from one phase to another.
Which of these applies to charged gases? The first equation is valid
when the subsidiary condition (6.24) holds: when it is possible to
change the volume of the individual phases at the expense of one
another, while keeping the total volume of the system constant.
Obviously this is not feasible in the system we are now considering.
The expansion of any phase would push a part of it into a region of a
different electric potential or, in other words, out of the phase. On
the other hand, it is clear that the equilibrium will not be disturbed
if we imagine each phase surrounded by a rigid heatconducting
envelope, i.e. if we assume the subsidiary conditions underlying the
formulas (6.27) and (6.29). We must conclude, therefore, that a
system composed of charged gases in
an electric field possesses, in its state
of equilibrium, a temperature T uni
form for all its parts while the pressure
is not uniform but " local," changing
with the potential from phase to phase.
The remaining condition of equilib
rium (6.31) is derived from the analysis
of processes of transfer of matter, from
one phase to another and from one
component to another, while the tem
perature of the system T and all the
local pressures of the phases are kept
constant. It applies to a charged system
to the extent to which such processes
are possible in it. Let us consider, for simplicity, a chemically pure gas
(i.e. having molecules of one kind only) in equilibrium at the temper
ature r, regarding the small volumes r and r r (Fig. 47) as two of its
phases with respective electric potentials 12 and Q'. The equilibrium
will not be disturbed if we imagine the small spherical surfaces L and
L' within the phases hardened and transformed into rigid adiabatic
FIG. 47. Equilibrium of charg
ed gas in electric field.
XVII 107 ELECTRON AND ION CLOUDS 269
shells. If the spheres are taken so small that the electric potential
due to the gas within them is negligible compared with the rest, we may
imagine this part of the gas removed without prejudice to the equilib
rium. Of course, we could have selected the spherical hole in r larger
by an infinitesimal amount and having the surface M (dotted line),
and that in r' slightly smaller with the surface M f . Let us adjust the
radii in such a way that the spherical layer between L and M contain
dN mols of the gas (in its original distribution), and that between
L 1 and M 1 an equal number dN' (i.e. dN = dN'), while the tem
perature and the local pressures remain the same. The same effect can
be obtained by expanding the sphere in r from L to M and contracting
that in r' from M 1 to L'. This last procedure provides the mechanism
we were looking for, a process for transferring dN mols of the gas from
the phase T' to the phase r without change of temperature or local
pressures in any part of the system. We can, therefore, apply the
formula (6.31) to it. Let the mol numbers and molal thermodynamic
potentials in r and r' be N,<p and N',<p', respectively. Changes of $
occur only in them, and the formula is reduced to <p dN + p'dN' = 0,
or because of the relation dN + dN' = 0,
<?  j. (17.04)
It is well to say here a word about the phase rule. We found in
section 43 that only three phases of a chemically pure substance can
coexist. However, that applies only to uncharged systems in which
temperature and pressure must be uniform throughout. On the
other hand, let us envisage a phases of the charged gas we are dis
cussing now. They are described by a + 1 thermodynamical variables
r, p (l \ . . . p (a \ and there exist between them a  1 relations (17.04).
Even if the electric potential differences are prescribed, two of the
variables can be chosen at random (no matter how large the number a
is), for instance, the temperature and the pressure in one of the phases.
If the electric potentials are not prescribed, there may exist still
further degrees of freedom.
The relation (17.04) is valid not only for a chemically pure gas: by
suitable use of semipermeable membranes it can be shown to hold for
any component of a gas mixture. However, we shall be interested
only in mixtures of perfect gases and, then, the following proposition
is obvious: if the relation
*/ = f r , (17.05)
holds for the component j, when it is pure, it will continue to hold,
when the component forms a part of a mixture. This follows from
270 TEXTBOOK OF THERMODYNAMICS XVII 107
the fact that perfect gases do not exercise forces on one another but
behave as if the others were not present.
In the particular case of a nondegenerate perfect gas, eqs. (17.03)
and (17.05) give
RT log W,/pd = ~ /(&' 0), (17.06)
(17.07,
This special .form of the relation could have been obtained also in
simpler ways, for instance, from Boltzmann's principle (4.69). How
ever, we shall have to fall back upon the general form (17.05) in treating
degenerate gases (section 111).
The other type of processes included in the equation of equilibrium
(6.31) are chemical reactions taking place within the individual phases.
In the interior of a phase the electric potential does not vary and merely
plays the role of an additional energy constant. The conditions are,
therefore, precisely the same in a charged gas as in an uncharged and
eq. (6.31) remains valid. In particular, when we consider the case of a
reaction of the type (6.46)
+ . . . + v0G0 = (17.08)
between perfect gases, the equation of equilibrium has the form (6.49)
of section 42 :
+ . . + r&t = 0. (17.09)
Comparing it with (17.05), we see that the equation is satisfied in
all the phases simultaneously so that we have equilibrium in every
part of the system. For nondegenerate perfect gases we have to use
for <f> the expression (17.03) obtaining the following explicit form of the
equilibrium condition
log pi = logK p  (Q/RT) },*&. (17.10)
K p is here exactly the same equilibrium constant as that defined by
eq. (8.15) in the theory of ordinary gas mixtures (section 51). The
sum Sv// = nj&vfij (where e$ is the charge on a molecule of the
component j) represents the change of electric charge in the reaction.
However, there is the law of conservation of electric charge, which
cannot be created or destroyed, so that the sum S^, = 0. The law
of equilibrium has, therefore, precisely the same form (8.17) in a mixture
XVII 109 ELECTRON AND ION CLOtf>3
of charged as of uncharged gases and can also be expressed in the
equivalent formulations (8.18) and (9.60).
108. Remark on heavy gases in a gravitational field. The con
siderations of the preceding section can be transferred in their entirety
to a system of heavy gases in a Newtonian gravitational field produced,
partly, by external causes, partly, by their own masses. The only
necessary change is to interpret 8 as a gravitational potential and to
write, instead of /, = n A e it the molecular weights M? = n A Mj, (Mj being
the masses of the molecules). The analogue to (17.06) is then
fifa  exp [M/W Q.)/RT\, (17.11)
which is simply a slight generalization of the wellknown barometric
law of Laplace's. In a gravitational field the equation of chemical
equilibrium takes the form
log pt = log K p  (0,/RT) T\/M/. ( 17  12 >
i
In most cases the sum S *>//*/, representing the change of mass in
the reaction, may be taken as zero. However, the law of conservation
of mass does not hold for matter alone, and there are phenomena for
which this sum is significant and may not be neglected (compare
section 133).
109. Thermoelectric potential differences in the classical electron
theory. As stated in section 106, the theory of electrons in metals
achieved some signal successes by assuming
that the free electrons form a perfect gas. In ^^^4^3 <2> I
the older (classical) theory this gas was re j
garded as nondegenerate, with an osmotic FIG. 48. Thermoelectric
pressure obeying the law (17.02) and eq. couple.
(17.07) resulting from that law. As an
application of these ideas, let us consider two bars (1) and (2), of two
different metals (Fig. 48), in contact at the junction J and forming
an open chain. Suppose that, in their electrically neutral state at
the temperature T, the two metals contain, respectively, z\ and $2
free electrons per unit volume. At the moment the two bars are
brought in contact, there exists on both sides of the junction a
difference of concentrations and, consequently, of osmotic pressures
of the free electrons. As we know from section 65, the osmotic pres
sure tends to equalize the numbers z\ and 212 and drives the electrons
from (1) to (2), if z\ > z 2 . How long will this socalled thermoelectric
272 TEXTBOOK OF THERMODYNAMICS XVII 109
action continue? As the electrons are driven toward bar (2), there
is set up an electric field: the potential of this bar will fall
(because of the negative electronic charge) and that of bar (1) will
rise. Equilibrium will be reached when the forces of the electric
field just compensate the force of the thermoelectric (osmotic) action.
Supposing that the system retains the uniform temperature !T, we can
calculate, for the state of equilibrium, the difference of electric poten
tials Oi fa between the two bars. We have only to apply the for
mula (17.07) of section 108, taking into account that pi/pz = zi/Z2i
according to (1.15), and R/f =  R/F =  k/e l
kT z\
$21 = 81 fo = log (17.13)
e 22
The assumption underlying this formula is as follows: the electron
density z of each metal is completely determined by its nature and,
possibly, by the temperature T, while the loss or gain of electrons
necessary for charging up the system is entirely negligible at a small
distance from the junction. In the immediate vicinity of the junction,
metal (1) carries a positive charge, through loss of electrons, and metal
(2) a negative one, through excess of them. The two opposite charges
form an electric double layer accounting for the drop of potential. 2
As long as the whole system is kept at the same temperature, the
thermoelectric potentials are not accessible to measurement. In fact,
their experimental determination would require a pair of wires going
from the free ends of bars (1) and (2) to a suitable measuring instru
ment. If the two wires are of the same material (3) we can denote the
thermoelectric differences with respect to the two bars by 831 and ^32.
We see, then, from the expression (17.13) that
832 = Osi + Oi2, (17.14)
so that both wires are at the same potential. It is obvious that taking
the wires of different materials will not remove the difficulty.
1 See footnote on p. 267.
1 Compare, however, footnote on p. 274. It was pointed out by some authors
that the energy necessary to form the double layer must be taken into account in
setting up the conditions of equilibrium (see: P. W. Bridgman, Phys. Rev. 27, p. 173,
1926). Theoretically this is true, but the terms arising from this cause are very small
numerically and share with the capillary surface effects the property of decreasing
in importance as the size of the system is increased (compare sections 39 and 81).
Since capillary forces are neglected in this and the following chapters, consistency
demands that the electric surface effects be treated on the same footing.
XVII 110 ELECTRON AND ION CLOUDS 273
In the present connection, we are interested in the thermoelectric
actions only in so far as they offer an interesting example of potential
differences in electron clouds. We shall see in section 139 that it can
be indirectly measured from the electromotive force of a thermo
couple with two junctions kept at different temperatures. Without
any doubt, the above theory explains the gist of the thermoelectric
phenomena, but it is somewhat oversimplified. This appears from
the experimental fact that thermoelectricity is structure sensitive: a
slight distortion in the lattice structure of a conductor, due to mechan
ical strain, produces an appreciable change of its thermoelectric
properties. As the theory stands today, it cannot account for the
observations, quantitatively, but does give the correct order of
magnitude.
It must be borne in mind that the formulas of this section apply
(within the approximation of the theory) only to conductors in which
the free electrons can be considered as forming a nondegenerate gas.
As was shown in section 105, this situation prevails in semiconductors.
Experimentally, it is known only that such materials possess unusually
large thermoelectric potential differences (compare Table 56 in
section 139). This is unquestionably due to the wide range of the
numbers z occurring in them. The thermoelectric conditions in
metallic conductors will be considered in section 113.
110. Contact potentials and thermionic emission in the classical
theory. Let us consider a conducting plate in vacuo: at its surface
the density of free electrons changes abruptly. The forces of osmotic
pressure (mentioned in the preceding section) tend, therefore, to
drive the electrons out of the conductor into the empty space beyond,
producing there a socalled thermionic cloud. We call the potential
of the conductor fli, the number of electrons in it (per unit volume)
01, and the same quantities in the thermionic cloud at the surface of
the conductor fl'i and z'l, and we apply the formula (17.13)
flj G'i = log^ (17.15)
e z i
The number z\ is determined by the nature and physical state of
the conductor, but the density z'\ of the cloud may assume any value.
What condition, then, makes the problem definite and fixes the poten
tial difference Oi 8'i and, through it, the number z'i? The action
of the osmotic pressure in producing the outer electron cloud is limited
by the fact that a definite amount of work W\ is required in order to
liberate one electron from the conductor , which is performed by the electron
274 TEXTBOOK OF THERMODYNAMICS XVII 110
while it passes through the surface. Therefore, the left side of
eq. (17.15) is determined by the relation
Qi  fl'i = Wi/e, (17.16)
whence ^ = ^ ^ (n A Wi/RT)  zi exp (bi/T), (17.17)
where the quantity bl = n A w 1/R (17 . 18)
is called the thermionic work function. 1
It is generally accepted that the force resisting the loss of electrons
and pulling them back into the conductor is the image force. As known
from the theory of electricity, a perfectly conducting plane acts upon
a point charge with a force equivalent to the attraction of the reflected
image of the charge. If the distance of the point from the plane is x
(its charge being e), the distance to its reflected image becomes 2x
and the image force is e?/4x 2 . The work necessary to remove the
electron from the position x = a to x = oo is, then, e?/4a. The actual
surface of the conductor is, of course, imperfect since it is constituted
of atoms with interstices through which the electrons can pass into the
outer space. Nevertheless, the above expressions for the image force
and the work done against it remain approximately valid as long as
x (or a) is not too small. To obtain the total work W\ that must be
supplied to an electron in order to get it free, one must attribute to a
a value of the order of the atomic distances (10~ 8 cm). The equations
Wi = e 2 /4a, fli  fl'i e/4a (17.19)
give then, in fact, the correct order of magnitude for the potential
difference, namely, a few volts.
Sometimes the question is asked how the electron cloud builds up
this difference of potentials. This question is illegitimate and based
on a misconception resulting from the fact that the term " potential
difference " is used in the above formulas in a sense slightly dif
ferent from that given it in most textbooks on electricity. 2 As under
1 Not to be confused with the "work function" denned in section 34. " n
1 The difference of electric potentials fij Q! between the points (2) and (1)
is determined by the work W n which must be supplied to a test body (of the charge e)
in order to move it from (1) to (2). Two different definitions are possible: either,
8a  81  lim (Wi2/e), when e approaches 0; or fti fli = Wu/e, when e is
finite. Most textbooks on electricity do not distinguish between the two definitions.
The second, obviously, takes in the work of any field induced by the test body and is,
therefore, unsuited for many theoretical purposes. However, it is the definition
which must be used in connection with Boltzmann's principle and thermodynamics,
since Cpot in eq. (4.69) is the energy imparted to a particle under the actual conditions.
In this sense (17.19) represents the actual difference of potential between the con
ductor and the outer space.
XVII 110 ELECTRON AND ION CLOUDS 275
stood here, the potential difference (17.19) exists from the start and is
not, to any material extent, built up by the electrons leaving the
conductor: the potential added by the electron cloud (so called, space
charge potential) is usually so small as to be negligible.
If a second conducting plate is placed at a short distance from the
first facing it in vacuo, it will also set itself in equilibrium with the
electron cloud and through it with the first plate. Two potential
differences relating to the plates are of interest. On the one hand, the
difference fli ^2 between the potentials in the interiors of the two
conductors depends (according to the theory of section 107) only on
their internal properties. It has, therefore, the same expression as
when the plates are in direct contact, and is identical with the thermo
electric potential difference discussed in the preceding section. On
the other hand, there is the socalled mutual contact potential 8'i  Q' a
of the two substances. It is defined as the difference of potentials
between two points in the electronic cloud which are, respectively, close
to the surfaces of the two plates. The potential 12' i at the surface of
the first plate is given by the eq. (17.16), that at the second plate by
the analogous formula $2 ^'2 = W^/e. Subtracting the two equa
tions, we find
tti  fl' 2 = Oi  &2 ~ (Wi  W 2 )/e. (17.20)
Substituting for &i 82 and W\ W% from the formulas (17.13)
and (17.18), we obtain for the contact potential in the classical theory
ffi  / 2 =  * [b,  6 2  r log (21/22)]. (17.21)
e
The second term in the brackets is usually very small compared
with the first.
The work function b can be measured either photoelectrically or by
thermionic currents. If we let light of a high frequency v fall upon a
metal plate, it emits photoelectrons having a kinetic energy JE^m.
Einstein's photoelectric equation
hv = Ek* + W (17.22)
expresses the hypothesis that the energy of an absorbed photon hv
is imparted to the photoelectron. The part W of it is spent in getting
through the surface of the metal, the remaining part appears as
kinetic energy. As the incident frequency v is decreased, Ewn becomes
smaller and smaller and vanishes altogether for a certain limiting value
hv<>  W. (17.23)
276 TEXlfcOOtf OP THERMODYNAMICS XVII 110
Light of lower frequency cannot produce any photoelectric effect,
and this permits the experimental determination of W. If we identify
W'with Wioi eq.(17.16), we find from (17.20) for the contact potential 1
fi'i  o' a    (y i  w) + * T log fa/**). (17.24)
e e
The method of thermionic emission for measuring b is based on the
formula from the kinetic theory of gases
n = z (RT/2*p)* 9 (17.25)
giving the connection between the number z of molecules per unit vol
ume in a perfect gas and the number n of those which pass from one
side to the other of any plane per unit time and unit area, n means
the molecular weight of the gas, in our case that of the electron
Me = 1/1821. This expression holds also for the electrons going in
either direction at the surface of the metal. However, of those going
outward the number rn are reflected at the surface, if r is the coefficient
of reflection. Only the remaining (1 r)n electrons come out of the
metal itself. The thermionic current is observed by making the heated
metal plate the cathode of a vacuum tube and applying a strong field
between the cathode and the anode. The electron cloud in the vacuum
is swept away by the field, but the conditions within the metal are not
appreciably changed and the electrons leaving it are still the same in
number, carrying with them the current
J = e(l  r), (17.26)
or substituting from (17.25) and (17.17)
J = A'T"e" r , (17.27)
where
A'  (R/2*n)*zie(l  r).
This is Richardson's formula for the thermionic current. It was
deduced by its author 2 from kinetic considerations and by H. A.
Wilson 3 from thermodynamics. For the reasons pointed out at the
end of the preceding section its validity cannot be claimed for metals
but is restricted to semiconductors.
1 It can be made plausible by statistical reasoning that Wi W is, in fact, the cor
rect identification. The data on semiconductors are not precise enough to test the
accuracy of formula (17.24).
* See footnote on p. 266.
See footnote on p. 266.
XVII 111 ELECTRON AND ION CLOUDS 277
111. The degenerate electron gas. We have seen in section 105
that the free electrons in metals must be regarded as a completely
degenerate gas. This fact was first pointed out by Pauli, 1 while
extensive study of metallic electrons, on the basis of Fermi's statistics,
is due to Sommerfeld. 2 The formulas of Chapter XVI, however, need
a small correction before they can be applied to electrons. It is caused
by the electron spins which obey the following rule: the spins of any
two electrons in the cloud are either parallel or antiparallel. We have,
therefore, two types of electrons with respect to the spin which we
shall call, for short, the first kind and the second kind. The statistical
weight and the intrinsic energy of either kind of electron are the same,
so that our gas is a mixture of equal parts of two components. Accord
ing to the remarks made at the end of section 105, the formulas of
Chapter XVI apply to each component separately, and it is easy to
infer from them the equations for the mixture. The specific volume
of each component is 2v, twice the specific volume of the mixture,
while the partial pressure of each kind of electrons is p/2, onehalf
of the total pressure. All we have to do is to substitute in all equations
2v, instead of v, and p/2, instead of p. As this leaves the product pv
unchanged, the only correction which the laws of degenerate gases
require in the case of electrons is a redefinition of the quantity 9
(replacing v by 2v)
instead of (16.02), where /u is the atomic weight of the electron
/LI, = 1/1821. Substituting the numerical values of the constants
e = 0.403 x io 5 v*r. (17.29)
There is reason to assume that in monovalent metals there are as
many free electrons as atoms. The molal volume of the free electrons
coincides then with that of the metallic material itself. For instance
itisv = 10.26 cm 3 mol^in the case of silver, giving = 1.91 X 10~ 6 r.
In other metals the values of 6 are of the same order of magnitude and
must be considered as small even at the highest temperatures (about
2300 C) accessible to thermionic measurement. It is, therefore, per
missible to use the expansion (16.10) and to obtain from (16.01)
W. Pauli, Zs. Physik 41, p. 81, 1927.
* A. Sommerfeld, Zs. Physik 47, pp. 1, 43, 1928.
278 TEXTBOOK OF THERMODYNAMICS XVII 111
When O 2 is negligible, the gas is completely degenerate. Its pressure
then becomes independent of temperature, being determined by the
specific volume alone which, in turn, is fixed by the nature of the
metal. We shall introduce the notation
RT
where !2< is called the inner potential. Hence its numerical value is
!2< =  25.85 tT H electronvolt. (17.32)
For instance, in the case of silver Q = 5.48 e.v. With the
abbreviation (17.31), eq. (17.30) takes the form
while the expression for the energy follows from (16.15) and (17.01):
 %pv +/Q + UQ. The entropy is determined by (16.17) and
becomes for small values of 9, according to (16.10),
2 / \2 T
'7 5
C1U3)
while the molal heat has the expression
X
Finally, the thermodynamic potential is expressed by
IT 2
^ +0 ~ + *' (17  35)
The electronic spin has an important bearing also on the entropy
expression for high temperatures. In of eq. (16.20), 2v must be sub
stituted for v, and this is equivalent to adding the term R log 2 to
the entropy constant, or log 2 to the chemical constant. The entropy
(16.21) and the thermodynamic potential (17.03) of the nondegenerate
electron gas retain, therefore, the form of their expressions, but the
constant JQ has an increased value. The chemical constant (14.15)
of the electron gas becomes (with /* = 1/1821):
jio, oi =6.180 (17.36)
or jio 0.826  1.
A similar increase of the chemical constant occurs also in the case
of gaseous atoms having an angular momentum and can be calculated
XVII 112 ELECTRON AND ION CLOUDS 279
in a quite analogous way. It will be better, however, to postpone this
calculation until section 120, where it will be treated together with the
chemical constant of diatomic and polyatomic gases.
112. Thermionic emission and thermoelectricity from the modern
point of view. We are now going to apply the results of the preceding
section to the equilibrium of a metal with the thermionic cloud outside
it. While the free electrons within the metal are in a degenerate state,
those forming the outer cloud are completely nondegenerate. In
fact, after deriving the conditions of equilibrium, under this assump
tion, we shall justify it by showing that the density of the cloud is
extremely low and the value of the quantity 9 in it very high. We
substitute into the condition (17.04), <p = <p f , the expressions (17.03)
for the outer cloud, and (17.35) for the metal. Noticing that
c p *  R, we find (with / =  en A )
log P = j + f !og r b/T, (17.360
with the abbreviation
 r _2 1.2121
(17.37)
where the potentials 0' and ft refer to the outer cloud and to the
metal, respectively. Neglecting second order terms,
b   (Q, +  a') (17.37')
K
With the help of (1.15) we throw (17.360 into the form
z  7^ exp (j  b/T), (17.38)
and obtain from (17.25) and (17.26) the thermionic current
 r)7V" i/p ,
(17.39)
A e (n A /2wpk)* exp j
120.2 amp cm~ 2 deg~ 2 .
This formula was first derived by Dushman l from a semiclassical
hypothesis of thermionic emission. Its derivation in terms of the
degenerate electron gas was due to R. H. Fowler. 2
The work function b receives, according to eq. (17.370, an interpre
tation different from that of section 110. The work which the electron
1 S. Dushman, Phys. Rev. 20, p. 109, 1922; 21, p. 623, 1923.
* R. H. Fowler, Proc. Roy. Soc. (A) 122, p. 36, 1929,
280
TEXTBOOK OF THERMODYNAMICS
XVII 112
has to do in getting through the metal surface is still determined by
the image force w which, in turn, fixes the outer potential difference
Q 12' as in eq. (17.17). However, the electrons bring with them the
energy eQ* depending on the inner potential (which has the nature
of a quantized kinetic energy, as is shown in the statistical version of
the theory). Only the difference bk = e(Q 0' + 0<) has to be
supplied by thermokinetic or photoelectric forces. The connection of
the thermionic work function with the photoelectric threshold fre
quency remains, therefore the same as in the classical theory, namely,
hvo  Kb.
Our knowledge of the inner potential fi, is not exclusively theoreti
cal: its existence and order of magnitude can be inferred from other
phenomena, for instance, from experiments on the diffraction of
electron beams falling on the surface of the metal. The values of ft<
found by such measurements are not at all accurate, as they are
affected by an uncertainty of d= 5 volts, but theoretical results obtained
from formula (17.32) do not lie far outside these limits. A few nu
merical data are given in Table 36. 1
TABLE 36
Photoelectric
Thermionic
Metal
kb/e
t C
kb/e
(calc.)
o o'
(calc.)
(obs.)
(volt)
If V
(volt)
Ag
4.73
20
4.08
925
5.48
10.2
18 5
4.56
600
Al
2.53.6
....
....
(5.57)
(8.6)
18 "
Au
4.82
20
4.32
1050
5.50
10.3
17 "
<
4.73
740
Cu
4.14.5
....
4.38
....
7.00
11.3
17 "
Fe
4.72
Hg
4.53
20
Ni
5.01
(7.35)
(12.4)
18 "
Pb
3.54.1
(3.9)
(7.9)
11 "
Pd
4.96
....
4.99
Pt
6.30
....
6.27
Rh
4.57
20
4.58
41
4.93
240
W
4.58
4.52
1 Taken from Hughes and Du Bridge (Photoelectric Phenomena, pp. 76, 235,
New York, 1932) and slightly revised. A more complete list will be found in the
article by J. A. Becker, Rev. of Mod. Physics 7, p. 123, 1935.
XV11 112 ELECTRON AMD ION CLOUDS 281
The inner potentials Q< in Table 36 are calculated on the assump
tion that there is one free electron per atom. This is sound in the case
of metals of the first group (Ag, Au, Cu) but less certain in other cases.
From the tabulated values of b one can determine z by means of
eq. (17.38) and find the molal volume and the quantity 9 in the ther
mionic cloud at different temperatures. For instance, in the case of
tungsten (neglecting the change of b with temperature) these data are
as follows :
TABLE 37
/C. 20 300* 600 1000 1500 2000 2300 3000 3800
v.... 1.1X10" 4.6X10" 3.4X10* 5.2X10" 1.6X10" 3.4X10" 2.5X10" 4.8X10* 5.1X10
6... 6.0X1078 7.6X10i 1.9X10" 7.8X10" 1.0X10* 2.1X10* 4.3X10* 3.7X10* 1.0X10*
We see from this table that 9 has exceedingly high values in the
electron gas outside the metal, so that the assumption of its being
nondegenerate is entirely justified.
In comparing Dushman'* formula (17.39) with experiments, it
must be remembered that it has its theoretical limitations. The
thermionic work function is not strictly a constant but has a slight
temperature dependence. This is due, in the first place, to the
neglected terms of the expression (17.35) and to the thermal expansion
of the metal which changes the molal volume v in it and also may
affect the image force, and through it, the potential difference 8 0'.
In the second place, the picture of electrons being completely free is
only an approximation; in reality, there are some mutual forces
between them and the metal ions which also could account for small
deviations from the constancy of b. Finally, the factor (1 r), which
is supposed 1 to be of the magnitude 0.9S, is again slightly temperature
dependent. On the other hand, the measurement of thermionic cur
rents involves an extremely difficult experimental technique whose
accuracy cannot be pushed beyond a certain limit. In view of this
situation, the agreement must be considered as quite satisfactory.
The results for A, as obtained experimentally, 2 do not agree with the
theoretical value 120.2, as shown in Table 38. This is probably due
to the work function b depending on T and containing a linear term
(b = b' + aT.) Therefore, the formula (17.39) can be rewritten in the
same form but with the new constants V = b aT, A' A exp a.
From the analysis of the measurements one obtains directly V, A' and
not 6, A, but, knowing the theoretical value of A, one can calculate a
and the true work function b.
L. Nordheim, Proc. Roy. Soc. (A) 117, p. 626, 1928.
L. A. DuBridge, Proc. Nat. Acad. 14, p. 788, 1928.
282
TEXTBOOK OF THERMODYNAMICS
TABLE 38*
XVII 112
Material
Authority
kb'/e
(in volt)
kb/e
A'/A
0X 10*
W
Bu Bridge
4.61
5.53
0.99 X 10
4.56
ThW
i <
4.61
2.63
1.05 X 10'
1.76
Pt
1 1
5.08
6.27
0.57 X 10*
3.41
K
1 1
0.645
2.21
1.37 X 10"
24.70
Cu (liqu)
4.55
3.54
1.65 X 10~ 4
6.91
Zr, Hf
Zwikker
3.85
4.84
0.57 X 10 a
4.08
BaO
Detels
2.48
1.47
1.89 X 10*
6.79
* See: S. Dushman, Rev. Modern Physics, 2, p. 381, 1930.
A curious and not entirely explained result of DuBridge's is that,
in the same metal but for different conditions of its surface, the
observed log A' is a linear function of the observed V (i.e.
log A' = &b' + const). The coefficients are listed in Table 38.
With respect to thermoelectric forces between two metals (1) and (2),
the qualitative considerations of section 4 remain valid. The equi
librium of the electrons in the two metals is determined by the same
condition <pi = ^2 The only difference is that we have to substitute
for the thermodynamic potential in either metal the expression (17.35)
since we consider both clouds as degenerate. Thus the potential dif
ference becomes
V 2 k 2 I 1
 2  (Q  Da)  T 2 (  f
Iz e? \Sm S
< 1M >
The expression fli $2<2 is independent of the temperature. It
represents a potential difference existing between the two metals under
all circumstances. It cancels out of every closed circuit and cannot be
observed by galvanometric or electrometric methods. Therefore, only
the second term need to be taken into account for the thermoelectric
difference:
821  0.611
x losr 2 ( f 
a
Q<2
volt,
(17.41)
where Qi, 12<2 are expressed in volts. This formula was first given by
Sommerfeld. 1 For the reasons outlined in section 109, we can hardly
expect from it more than the correct order of magnitude. Experi
1 A. Sommerfeld, note on p. 277. This formula does not represent the e.m.f.
in a thermoelectric circuit, which is given at the end of section 139.
XVII 113 ELECTRON AND ION CLOUDS 283
mental data and a thermodynamical discussion of thermoelectricity
will be given in section 139.
As to the contact potential of metals, it can be obtained directly
from the formula (17.37), substituting for 12i Ife the expression
(17.40),
This relation was tested by Millikan, Lukirsky and Prilezaev, and
Olpin * and was confirmed in the case of metals with very clean
surfaces.
113. lonization in the solar atmosphere. In their systematic
study of solar radiation the astronomers succeeded in analyzing spec
troscopically the light coming from the different altitude layers of the
chromosphere. They found that the emission of the lower levels mostly
consists of arc lines (characteristic of the neutral atoms), while the
spark lines (belonging to the ionized atoms) appear higher up and
grow in intensity with altitude at the expense of the spark lines.
This result seemed puzzling because the lower layers of the sun are
hotter and, in the laboratory, the spark lines are enhanced by raising
the temperature. The reason for this discrepancy was explained by
M. N. Saha, 2 who pointed out the influence of the pressure on the
degree of ionization and applied the thermodynamical equations of
chemical equilibrium to the problem. There are charged particles of
both signs in the solar chromosphere, but at the temperatures and
pressures prevailing there the phenomenon of pairing (mentioned in
section 106) is quite negligible. Therefore, the form of the equations
of equilibrium is not appreciably changed either by the charges on the
electrons and ions or by the gravitational field of the sun (section 108).
These conditions also exclude degeneration so that we can apply the
equation of chemical equilibrium (14.16) for classical perfect gases.
The reactions which we have to consider as taking place are those
of the single and double ionization of atoms by splitting off one or two
electrons. In the symbols of section 42
#++GLG n = and C++ + 2GL  G n  0, (17.42)
where" 6+ and C++ refer to the positive ions, GL to the electron and
G n to the neutral atom. The formula (14.16) needs a slight correction;
1 R. A. Millikan, Phys. Rev. 7, p. 355, 1916; P. Lukirsky and S. Prilezaev,
Za. Physik 49, p. 236, 1928; Olpin, Phys. Rev. 36, p. 251, 1930. For a discussion
from the point of view of the statistical theory see: Mitchell, Proc. Roy. Soc. (A)
146, p. 442, 1934; 153, p. 513, 1936.
* M. N. Saha, Phil. Mag. 40, pp. 472, 809, 1920.
284
TEXTBOOK OF THERMODYNAMICS
XVII 113
there must be added log 2 to the chemical constant of the electron
(section 111). However, this is not all : it will be shown in section 120
that a similar term log g belongs to the chemical constants of the
atoms and ions (because of the statistical weight of the angular
momenta of these particles. Since the atomic weights of the atom
and the ion are practically the same, we can write f S> A logio MA =
f v logio M ~ 4.892v, where p e means the atomic weight of the elec
tron. For monatomic gases c p = %R, and the heat of reaction be
comes, according to (3.28), Q = 2^x* = At/o + fyRT. If we con
tinue to measure the pressure in atmospheres but take as the unit of
heat of reaction not the erg but the electronvolt (1 erg/mol =
0.9646 X 10 12 e.v.), the equations of equilibrium become for the
two reactions (17.42)
log (P+P/P*)  ~ 6.180 + logio
 f ^
f logio T
 12.360 +
^
+ 5 logio T
5037 AC/o
5037 AC/o
(17.43)
The values of AJ7o and g for the most important ingredients of the
solar atmosphere are given in Table 39. The data of this table refer to
the normal states of the particles', the statistical weights of the nuclear
spins are not included in g because they are the same in the neutral
atom and the ion and cancel out of the reaction.
TABLE 39
Particle
I
Al/o
(in e.v.)
Particle g ( .^ v }
Particle
g
(in e.v.)
H
H +
1}
13.53
Ca 1
Ca + 2
6.09
11.82
Zn
Zn +
i!
9 36
17.86
He
Na
Na+
1}
24.46
5.12
Sr** 1
Sr + 2
Sr ++ 1
5.67
10.98
Cd**
Cd +
1 J
II
8.96
16.84
Rb
Rb+
1}
4.16
Ba 1
Ba+ 2
5.19
9.96
Hg
**g+
ii
10.38
18.67
Cs
?}
3.87
Ba++ 1
rig +4.
1'
c...
2!
23.4
XVII 114 ELECTRON AND ION CLOUDS 285
We see from eqs. (17.43) that the ratio P+/p not only is an increas
ing function of the temperature but also depends on the pressure of
the free electrons, being inversely proportional to it. This pressure is
an appreciable fraction of the total pressure and rapidly decreases with
the general attenuation of the chromosphere as the level is raised. The
equation, therefore, gives a qualitative explanation of the observations
described above. Moreover, it accounts for the observed fact that,
the lower the ionization potential, the lower also the level at which
the appearance of the spark lines sets in. On the other hand, it is
very difficult to calculate p+/p n quantitatively because the pressure
of the electrons />_ depends on many factors. All the elements of the
mixture which constitutes the solar atmosphere contribute to it, but
even if we had atoms of one kind only the degree of ionization would
depend on the excited states and its computation would be a difficult
problem. 1 However, if our object is only to obtain a check of the
thermodynamical formulas (17.43), we can get around this difficulty
by choosing one of the elements of the solar chromosphere as a stand
ard. Suppose we have measured, for different levels of the sun, the
intensities of the arc lines and the spark line of this standard substance
and inferred from these measurements the relative abundance
/>+<> //><> of its ions and neutral atoms. Writing eq. (17.43) for the
standard element and for any other element, we eliminate /> and
obtain
The theory was tested by similar devices and received a fair con
firmation, but it was pointed out by H. N. Russel 2 that great accuracy
in the agreement cannot be expected. The gases of the solar chromo
sphere are not strictly in a condition of thermodynamical equilibrium
since they are exposed to the onesided radiation coming from the
lower and hotter regions of the photosphere.
114. Ion concentrations in electrolytes. Galvanic cells. The
theory of dilute solutions also offers an important field of applications
for the results of sections 106 and 107. A large class of solutes, the
electrolytes, undergo a partial or complete dissociation into electrically
charged components (ions). This is particularly characteristic of
acids and salts in aqueous solutions: when their molecules dissociate,
1 It was treated by Milne, R. H. Fowler, and C. G. Darwin. The most compre
hensive review of the whole field of ionization in the solar atmosphere is due to
A. Pannekoek (Handbuch der Astrophysik, Vol. Ill 1, pp. 257350. Berlin 1930).
* H. N. Russel, Astrophys. J. 79, p. 317, 1934.
286 TEXTBOOK OF THERMODYNAMICS XVII 114
the hydrogen or the metal acquires a (single or multiple) positive
charge and is called the cathion, or positive ion, the acid radical gets
a negative charge and constitutes the anion or negative ion. As an
example we take cadmium sulfate which, dissolved in water, dissoci
ates according to the formula
Cd++ + SO 4 '  CdSO 4 = 0,
where the superscript ++ indicates a bivalent positive ion (i.e. one
carrying two elementary charges).
We know from section 68 that nonelectrolytic dilute solutions
behave like perfect gases. If the conditions of equilibrium are not
changed by the electric charges of the ions, we should expect the laws
derived in that section, especially the mass law (9.60), to apply also
to electrolytic solutions. For instance, the above reaction would give
X++X /Xn = K(T),
where x++, x , x n are the respective mol fractions of the positive and
negative ions and of the neutral (undissociated) salt. In the more
general case, when the molecule is ionized into vi, i>2, ... V T positive
and negative ions, their total number being v t ~ v\ + . . . + v rt the
equation takes the form
vid + P2G* + ... + v T G r  G n = (17.45)
and the mass law has the expression
X) ^ log Xj  log x n = log K. (17.46)
i
This equation is of the same type as those considered in section 52,
and we can draw from it the conclusion that a certain fraction of the
salt molecules is ionized (dissociated into ions). In fact, experiments
show that the equilibrium constant K is fairly large in these reactions,
so that solutions of low concentration are almost completely ionized.
The hypothesis that dilute electrolytes are completely dissociated
into ions was first advanced by Arrhenius, who adduced in its support
their high osmotic pressures and low freezing points. Moreover, he
showed that the electric conductivity is in agreement with this assump
tion, being nearly proportional to the molality of the dissolved salt.
The work of Arrhenius and his successors had a great influence on the
development of thermodynamics, on one hand, and laid the founda
tions of the science of electrochemistry, on the other. The laws of
electrolytic conduction, as developed by Kohlrausch, afforded a means
XVII 114 ELECTRON AND ION CLOUDS 287
of determining the mol fractions Xj of the individual ions and gave a
rough confirmation of the formula (17.46). However, as the technique
of these measurements grew more and more accurate, it was found
that the law (17.46) is rigorously valid only for a limited class of sub
stances, the socalled weak electrolytes. To this group belong a number
of organic acids and their salts. On the other hand, the common
inorganic acids and their salts form the class of strong electrolytes which
exhibit appreciable deviations from the formula (17.46) even in very
low concentrations.
At first, it was doubtful whether these discrepancies were due to
the inadequacy of the formula itself or to the failure of the conduction
measurements to give the correct mol fractions. The investigations
based on the freezing method, described in section 78, were under
taken to decide this question. They led to the result that eq. (17.46),
actually, is not valid for strong electrolytes, or in other words, the ionic
clouds of strong electrolytes do not obey the laws of perfect gases.
As was shown in section 77, the form of the mass law can always be
retained if the numbers Xj are replaced by the effective mol fractions
or activities a,:
JjiV log a, log a n = log K. (17.47)
Putting for the neutral salt a n = x n and making use of the notations
of section 79, especially of the activity function 7 defined by (11.32),
we can write (for dilute solutions)
vj log Xi log x n = log K v, log 7. (17.48)
The data from conduction measurements turn out to be in excellent
agreement with those obtained by the freezing method: when the
proper values of 7 are substituted, the formula (17.48) is confirmed with
great accuracy. The theoretical reasons for the abnormal behavior
of strong electrolytes will be discussed in the next section.
The notion of electrolytic ionization affords an insight into the
mechanism producing the e.m.f. (electromotive force) of galvanic cells
which were treated in section 37 from the more formal point of view
of energy relations. In all the constructions mentioned there the
chemical reaction is between two metallic salts, MR and M'R (with
the same radical R) which ionize in aqueous solutions according to
equations similar to
MR  M + + , M'R = M'+ + R, (17.49)
and in which the reaction M'R + M > MR + M' is exothermic.
288
TEXTBOOK OF THERMODYNAMICS
XVII 114
The current is, in part, due to the negative radical ion R traveling
through the cell from the vicinity of the positive to the negative plate. 1
We are now able to say what force drives the ions and produces the
current: it is the force of osmotic pressure. In fact, the R_ ions react
at the negative plate with the metal M of which it consists. Therefore,
their concentration and partial osmotic pressure P_ sinks below the
value P'_ which it has at the positive plate. When the resistance of
the outer circuit is very large, an electric difference of potentials
fl' 12 builds up which almost completely compensates the e.m.f.
(electromotive force) of osmotic pressure.
According to (17.06) we have then (since/ = F)
(17.50)
  f log (P'/P),
provided the cloud of R~ ions in the solution can be regarded as a
perfect gas. The explanation for the movement of the M'+ ions to
the positive plate is also osmosis (since in equilibrium the total osmotic
pressure must be everywhere the same), but the calculation of their
concentrations would lead us too far into electrochemistry.
It is essential for the operation of the cells described in section 37
that its electrodes consist of two different metals; they are, therefore,
called electrode cells. There exist cells of a different kind, known as
concentration cells, an example of a modern concentration cell is as
follows: 2
Ag
AgCl
(solid)
KG
solution (m)
KHg
amalgam
KG
solution (m f )
AgCl
(solid)
Ag
The first half of this chain (from Ag to the amalgam KHg) repre
sents an electrode cell with the reaction AgCl + K = KCl + Ag,
which is strongly exothermic. If this half chain were left to itself, the
molality m of KCl in the chamber filled with potassium chloride solu
tion would increase producing a current in the direction to the left
(i.e. from KHg to Ag). In fact, in this case, the K+ ions move to the
left and replace some of the silver in the AgCl paste (the reduced
silver being deposited on the Ag electrode). At the same time the
Cl_ ions move to the right and react with the potassium metal in the
amalgam. In both processes KCl is produced and added to the solu
tion. On the other hand, if an electric current is sent through the half
1 In the example of the Weston cell (Fig. 11, p. 95) the radical (R.) is SO 4 ,
the negative plate (M) consists of Hg and the positive (M') of Cd.
* Maclnnea and Parker. J. Am. Chem. Soc. 37, p. 1445, 1915.
XVII 114 ELECTRON AND ION CLOUDS 289
chain in the direction to the right all processes are reversed and the
molality of potassium chloride solution must decrease. Now, the
second half of the concentration cell is of the same construction and of
opposite orientation: the resultant current will be, therefore, the
difference of the currents produced in the two half chains and will flow
through both of them. If we assume m' > m (see below), it will have
the direction from KHg to Ag, in the first half, and from Ag to
KHg, in the second. Therefore, the molality in the first KC1 cham
ber will increase at the expense of that in the second. There is no cur
rent when m r = m : the driving power is due to the difference of osmotic
pressures in the two chambers.
All this is also characteristic of the general case when the reaction
is given by eq. (17.45). The essential feature of every concentration
cell are two chambers C and C filled with the same solution in different
concentrations. In the beginning, when the chambers are freshly
filled, the solutions are electrically neutral (as a whole, containing
equal positive and negative charges) so that the osmotic forces are
entirely uncompensated by electric potentials. Under the ordinary
conditions of construction, there is no tendency for potential differences
to build up, and the solutions can be regarded as neutral during the
whole length of their operation. Owing to these forces, the ions pass
from the chamber of higher osmotic pressure into that of lower in
numbers proportional to the coefficients v$ of eq. (17.45) as is implied
in the electric neutrality of the solutions. We shall calculate , the
e.m.f. of such a cell, only for the case of low concentrations when the
electrolyte (in both chambers) can be considered as completely ionized.
The change of thermodynamic potential which the ions undergo in
being transferred from the chamber C to the chamber C is then
(referred to 1 mol of the neutral electrolyte) A* = %v h fo\ ? h ).
The left side, according to (5.35), is the (negative) nonmechanical
work done by the system APT  W  EJ, where / is the total
charge of the ions pertaining to 1 mol. On the right side, we substi
tute the activities from formula (11.17)
JE   RT y>* log (aV**). (17.51)
If we denote the valency of the ion h by <r h (taking it positive or
negative, according to the sign of the charge)
290 TEXTBOOK OF THERMODYNAMICS XVII 11$
Further, making use of the abbreviation (11.32), we can write
E   (RT/J){v 9 log ( 7 '/7) +Z* log (x'k/tk)} (17.52)
The measurement of electromotive forces of concentration cells offers
a method for the determination of relative activity functions of electrolytes.
Numerical data for the activity function will be given in the next
section.
115. Strong electrolytes. All the results so far obtained in this
chapter rest on the following hypothesis made in section 106: the
mutual electric energy of the particles in a small volume r is propor
tional to the square of the number of particles in it and, therefore,
negligible. However, this argument is sound only as long as the
electric forces do not affect the uniformity of distribution of the
particles, and it ceases to be strictly applicable in the case of mixtures
of positive and negative ions. Owing to the mutual attraction, each
ion has the tendency to hover in the neighborhood of other ions of the
opposite sign. The mean distance between unlike ions is smaller than
that between like ones. This gives rise to energy terms, proportional
to the number of particles, which may become significant under
certain circumstances. It had been suspected for a long time that
these additional electric terms may account for the abnormalities in
strong electrolytes mentioned in sections 79 and 114. If this point
of view is accepted, the sharp line of division between dissociated and
undissociated molecules disappears: the distances between the ions
can assume all values and are distributed according to the laws of
probability. The transition from association to dissociation is gradual,
and all ions can be considered as free, in this sense.
The problem was formulated with complete clarity and partially
treated by Milner. 1 It was made better amenable to quantitative
evaluation by Debye and Hiickel, 2 who gave a simple method for the
approximate calculation of the effect at low concentrations.
Let us consider a neutral solution containing ions of several kinds,
the type (j) being represented by Ni mols. We denote their valency
by <r/, which is positive or negative, according to the sign of the charge
on this ion. The total charge in the solution is then
0, (17.53)
Milner, Phil. Mag. 23, p. 551, 1912; 25, p. 742, 1913.
P. Debye and E. HUckel, Phys. Zs. 24, pp. 185, 305, 1923.
XVII 115 ELECTRON AND ION CLOUDS 291
where F = en A is the Faraday constant or the charge per 1 mol of
monovalent ions. We denote farther by
ztNtnJV, (17.54)
the mean density, i.e. the number of ions of the type j per unit volume.
Let us now focus our attention on a " preferred ion " of the kind (1)
and let us ask how the other ions are distributed around it. The
answer is given by the Boltzmann principle (4.69). Suppose that, at
a point P, its electric potential is 81 ; in the time average, the density
of jions at this point will be C exp (e<rjQi/kT). Let the potential
81 be counted "from infinity", i.e. from a distant point where the
field of the preferred ion does not make itself felt (Qi = 0) and the
density has the normal value /; then the constant of proportionality
has the value C = */. The total density of electric charge at the point
P, produced by ions of all kinds, is, therefore,
exp ( eat Qi/kT). (17.55)
j
The potential d is produced by the preferred ion, in part directly,
in part indirectly through the action of the " atmosphere " of other
ions which hover about it. Assuming the validity of the Coulomb law
and, on the average, a spherically symmetrical distribution of the
atmosphere, fij. must satisfy Poisson's equation in the form
1
r 2 dr \ dr / D
where D is the dielectric constant of the solvent. Eliminating p from
(17.55) and (17.56), we obtain a differential equation for the determina
tion of QI. We restrict ourselves to the case when the exponent
<TJ e$li/kT is so small that its square can be neglected. Because of
(17.53), the term of zero order vanishes, leaving
2 m ,
H  = *r 81, (17.57)
r dr
*%*%* <" 58 >
The integral of eq. (17.57) is
! = ". (17.59)
292 TEXTBOOK OF THERMODYNAMICS XVII 115
The coefficient A\ must be determined from the "boundary
conditions." Let a\ be the mean distance of nearest approach of the
other ions to the preferred ion: within the little sphere of the radius a\
the field is not influenced by the ionic atmosphere and the potential in
this region may be taken as obeying the simple Coulomb law
tf x  ? + ODI, (17.60)
while outside this region it follows the law (17.59). At the surface of
the sphere r = ai, there must be continuity of potential (&'i = fli)
and of dielectric displacement (dQ'i/dr = dtti/dr). This leads to
ffl e exp
Without the ionic cloud about the preferred ion, its potential
(against infinity) would be represented merely by the first term of the
expression (17.60). Therefore, QOI gives us the potential difference,
built up through mutual interaction, of the ion with respect to the mean
potential in the solution. In a similar way the other ions assume the
potentials
w TT (17  62)
D 1 + *&/
In arriving at this formula, a dual capacity was attributed to
every ion. On the one hand, it acted as a " preferred ion " collecting
an atmosphere around it; on the other, it formed part of the atmos
pheres of all the other ions. It is clear that it is not always permissible
to separate and superpose these two functions of the ion, as is done in
DebyeHuckel's theory. The superposition implies, ultimately, a
linearity of the equations controlling the electric field and comes back
to the same restriction which was made in eq. (17.57). It seems,
therefore, risky trying to improve the theory by taking into account
terms of higher order in the exponent of (17.56), as has been done by
some authors. 1 The question needs a closer investigation whether
the theory is not restricted by its very method to cases in which
(ffjeQi/kT) 2 can be neglected. Another idealization of the theory
lies in the schematic character of the " boundary conditions " which
lead to the formula (17.62). In view of them, the dependence of the
potential Ooy on the ionic radius a, can hardly be considered as entirely
trustworthy.
1 Gronwall, LaMer, and Sandved, Phys. Zs. 29, p. 358, 1928.
XVII 115 ELECTRON AND ION CLOUDS 293
We shall show in section 138 that the change of internal energy of
the solution produced by the potentials (17.62) is a rather complicated
one. The work function SF, however, can be calculated without much
trouble. In order to do so, it will be well to bring out the thermo
dynamical significance of our results. An ion of the charge cr, 6 acquires
the potential flo/' let us consider an imaginary, ideal way of producing
such a distribution of charges and potentials. Suppose that e is not
constant but that the charge of every ion can be changed and built
up from to the normal value. 1 The work which must be done in
order to increase all the charges from <r,e to <r,( + de) (e being the
same for every ion) is known from the theory of electricity to be
DW* = Qo/ffi <fct where the summation is extended over all the
ions. Since there are n A N, ions of the kindj, DW 9 = n A / .ffjN&oide.
j
The total work of building up the distribution at constant temperature
and constant volume is, therefore,
W.   n A I y\yQ ,tf/ de. (17.63)
It was shown in section 36 that the nonmechanical work done in a
reversible process, at T = const, V = const, is oppositely equal to
the increase of the work function. This increase, due to the electric
charges and potentials of the ions, we denote by ^, = W 9 . The
expression (17.63) is somewhat simplified if we use, instead of the
individual ionic radii a/ of eq. (17.62), a mean radius a. In fact, the
accuracy of the theory is hardly sufficient to justify such distinctions
and, moreover, the term *a/ = *a in the denominator is altogether
neglected in our ultimate applications. According to the definition
(17.58) of K we can write, then,
(17.64)
The partial thermodynarnic potential can be obtained directly
from the work function by means of eq. (5.39)
_ _. kTV "* ( d "\ .
T 4ir l + Ka\dN h / VtT 2D 1 + *a " '
This is the increase of the thermodynamic potential due to the
electric fields of the ions. If we define the activity coefficient of the
ions h as log a h = ?eh/RT (section 78), this means that we compare the
1 R. H. Fowler, Statistical Mechanics, p. 318, Cambridge, 1927.
294 TEXTBOOK OF THERMODYNAMICS XVII 115
activities in strong electrolytes with the theoretical ones in dilute
solutions at the same volume, whereas experimentally they are compared
at the same pressure. This difference, however, is of slight significance
in view of the very small compressibility of water and other condensed
solvents. 1
In very dilute solutions K is comparatively small, and Ka in the
denominator can be neglected. From (17.58)
or in terms of molalities (V = 1000, AT, =
log a* =  ,^* H ( ^ > V 2 , ) (17.66)
In water at the temperature of 25 C the dielectric constant is
D = 78.8 leading to the expression
logio <* h =  0.356 fft?
if we denote by m the molality of the neutral electrolyte originally dis
solved (mj = Vjm). Hence the activity function, defined by eq. (11.32),
is given by
log
lo 7 . _
while the coefficient becomes 0.345 for the temperature C. This is
the socalled limiting law for very dilute electrolytes due to Debye and
Hiickel.
The agreement of the limiting law with the measured activities is
very good for ions of low valency as appears from Table 40.
The very accurate measurements by Neumann, 2 on silver chloride
and barium sulfate, show, however, that the limiting law is not rigorous
even for lowvalency salts. In plotting the measured y against m**
this author finds curious oscillations of the curves around the theo
retical value (17.67). The deviations, although very small, seem
outside the limit of experimental error. Highvalency electrolytes
1 The pressure increase due to the effect here considered is only a small fraction
of 1 mm Hg. According to (9.42), (11.35), and (17.67), in water of 25 C it is:
A  A  0.006v e m log y  0.12(2<r,yym)^ mm Hg.
E. W. Neumann, J. Am. Chem. Soc. 54, p. 2195, 1932; 55, p. 879,
XVII 115 ELECTRON AND ION CLOUDS 295
were investigated by La Mer and coworkers, 1 who found large dis
crepancies with the limiting law.
TABLE 40
ACTIVITY FUNCTIONS OF STRONG ELECTROLYTES
Type <TI = 1, <r a = 1, n 1, v* 1
0.005
0.932
0.929
0.930
0.926
 2
0.005
0.753
0.785
0.800
0.734
Type o\ = 2, 02 = 2, v\ 1, v\ 1
m .............. 0.001 0.002 0.003 0.005
7 ......... (theor) 0.746 0.661 0.601 0.519
BeSO 4 . . . . (exp) 0.754 0.670 ..... 0.534
CdSOi.... " 0.754 0.671 0.621 0.540
CaS 2 O 8 ... " 0.754 0.674 ..... 0.540
When the solutions are not very dilute (presumably, m > 0.005,
for low valency salts), the term KO, in (17.65) can no longer be neglected.
In fact, the measured activity functions usually fall, at the higher
concentrations, below the values of formula (17.67). However, in
the present state of the theory the dependence on a is not accurately
enough known to make its quantitative discussion worth while. 2
i La Mer and Mason, J. Am. Chem. Soc. 49, p. 410, 1927; La Mer and Cook,
ibidem 51, p. 2622, 1929; La Mer and Goldman, ibidem 51, p. 2632, 1929.
* A review of the work done on the theory of strong electrolytes by statistical
methods will be found in an article by O. Halpern (J. Chem. Physics 2, p. 85, 1934).
A critical discussion of the foundations of the theory was given by R. H. Fowler,
(loc. cU.)
m
0.001
0.002
0.003
. . (theor)
0.964
0.949
0.938
HC1 . . . .
.. (exp)
0.965
0.953
HNOs
4
0.953
AgNO 8 .
44
0.940
AgC10 3 .
KC1
41
41
. . .
0.940
m .
Type
01
0.001
= 2, a a = l,
0.002
v\ 1,
0.003
CaCl 2
..(theor)
(exo)
0.881
888
0.836
0.850
0.803
BaBra
44
0.829
H*SO*. .
44
0.876
0.825
CHAPTER XVIII
THEORY OF SPECIFIC HEATS
116. General considerations. We had occasion to give the expres
sions for the specific heats of a few ideal systems, such as perfect gases
(nondegenerate and degenerate) and the Einstein model of a solid.
Purely thermodynamical reasoning does not enable us to say much
about the heat capacities of the real substances of nature except that
they must vanish at T = 0. The internal energy of a system com
pletely determines its equation of state through eq. (13.22), but the
reverse is not true: the knowledge of the equation of state is not
sufficient for calculating the energy since eq. (16.12)
still contains an unknown function of the temperature which usually
gives the major contribution to the specific heat. However, heat
capacities are of such fundamental importance as material for the
applications of thermodynamics that a textbook of this branch of
science would be incomplete if it did not contain some information
on the status of our theoretical knowledge about them, even if this
knowledge is obtained by nonthermodynamical methods. It seems
appropriate, therefore, to include a chapter partially devoted to the
way in which the kinetic theory of matter approaches the problem of
specific heat.
The fundamental tool of the kinetic treatment is eq. (4.70) of the
Boltzmann principle. It is true that there are other types of statistics
which we have described in section 104. However, as it was stated
there, they apply only when many identical particles are distributed
over energy states determined largely by joint external conditions.
On the other hand, in such systems as an assembly of rotating mole
cules, each particle rotates around its own center of gravity and its
quantum states are determined by conditions peculiar to itself as if it
were alone and not part of a system. The same thing applies to the
other systems occurring in the theory of specific heats, for instance,
296
XVIII 116 THEORY OF SPECIFIC HEATS 297
linear oscillators. There is, therefore, little occasion to use the statistics
of Fermi or EinsteinBose, and the Boltzmann principle reigns supreme
as far as these investigations are concerned. If a particle can assume
the quantum states 0,1,2, ... with the respective energies eo, ei, 82, ...
the principle postulates that the number of particles in the state / be
Zi = Cexp ( Bi/kT). The sum of the numbers Zj represents, obvi
ously, the total number Z of particles in the system
Z = c exp (ci/*r)  CY. (18.02)
The expression Y = S exp ( i/kT) is usually called the sum of
states. Every quantum state of the particle must be represented in it by a
term. If the same energy level e< belongs to several (gi) different
quantum states, the corresponding term must be repeated g\ times.
It is, therefore, more convenient to write it in the form
V = Z ft exp ( ei/*r). (18.03)
We call the integer gi, as in section 111, the statistical weight of the
level /.
The energy of all the particles in the state / is
X exp ( ei/kT), and the total energy of the system
U = C]C gi*i exp ( ei/*r). (18.04)
The sum in this expression can be obtained by differentiating Y
with respect to (l/*r,) i.e. U =  CkdY/d(l/T). Eliminating C
with the help of (18.02)
If we refer our considerations to 1 mol of matter, the number of
particles is Z = n A (Avogadro's number n A = R/k), and the molal
internal energy becomes
u =  R a log F/a(i/r>. (IBM)
We may consider this as the fundamental formula of the statistical
theory of specific (molal) heats, since they can be derived from u by
the formula c = du/$T or
R a 2 log Y
< 18  07 >
298 TEXTBOOK OF THERMODYNAMICS XVIII 117
Comparing (18.06) with the relation (5.S6), we find as the thermo
dynamical interpretation of the sum of states
* =  RT log Y, (18.08)
where \l/ is the molal work function defined as ^ = u Ts. There
fore, the entropy has the expression
s ^Ar + *"* r < IM
When the temperature T is very low, all the terms of the sum of
states (18.03) become negligible compared with those of the lowest
energy eo. The sum is reduced to Y = goesp ( eo/Jfer), whence
RT log Y = n A o + RT log go and s = R log go It may seem,
at first sight, that this result is in contradiction with Planck's formula
tion (13.19) of Nernst's postulate. However, these formulas are no
longer valid in the vicinity of T = 0. The expression (18.03) refers
to the rotational and oscillatory motions of the molecules, and its use
implies the assumption that the probabilities of these motions can be
evaluated apart from the translational movements. This ceases to be
true under conditions of temperature and density in which the statis
tics of Fermi or Einstein Bose begin to play a role for the translational
degrees of freedom. We refer, in this connection, to the remarks made
in section 111 with respect to the entropy of the electron gas. In that
example the statistical weight is go = 2, as there exist two kinds of
spin. In the nondegenerate state the entropy contains, therefore, the
term R log 2 (entropy of mixing), but in the degenerate gas this term
disappears because the two kinds of electrons are regularly and
uniquely arranged over the quantum states.
117. Diatomic gases. The classical kinetic theory inevitably
leads to the principle of equipartition of energy (4.71) : each degree of
freedom takes up the kinetic energy %RT per mol. It b'assumed that
the molecule of a monatomic gas has only 3 (translational) degrees of
freedom, the molecule of a diatomic gas 5 (3 translational, 2 rota
tional). The corresponding molal heats are c v = $R and c p = J?,
respectively. If the gas does not strictly obey the equation pv = RT,
the resulting correction is taken care of by eq. (18.01) or by its deriva
tive with respect to T
/V^2jA
dv+f'Cn, (18.10)
while c p can be obtained from (4.27). The main success which the
classical theory has to its credit lies in the sphere of monatomic gases,
XVIII 117 THEORY OF SPECIFIC HEATS 299
whose specific heats it represents accurately. It is not so satisfactory
in application to diatomic substances (compare section 15). For the
more perfect permanent gases of nature the agreement is fair in the
region of room temperatures, but there are large deviations both for
low and very high temperatures.
The following treatment is based on the theorem of mechanics that
the kinetic energy of any system (molecule) can be represented as a
sum of the two terms : the kinetic energy of the center of gravity and
the energy of the motion within the system (molecule) relatively to the
center of gravity. Correspondingly, the internal energy (as also the
entropy and the specific heat) may be divided into two parts. It
happens, moreover, that the contributions of the rotations and vibra
tions (though not rigorously additive) are separable with a sufficient
degree of approximation 1
u = u t + u r + u vi . (18.11)
The first belongs to the translational motions and is calculated
by the classical formula
wo, (18.12)
since the center of gravity is completely determined by the three
coordinates of translation, assuming that degeneration of the type of
Chapter XVI is negligible (compare preceding section). The second
and third parts are due to the rotational motion of the molecule as a
whole or to the vibrations of the atoms inside it. They are computed
with the help of the quantum eqs. (18.03) and (18.06).
As to the rotational heat c r , a good approximation is often afforded
by the socalled dumbbell model. The diatomic molecule is considered
as consisting of two material points at an invariable distance from each
other rotating around their center of gravity in a joint plane. Quantum
dynamics gives for the energy levels of the dumbbell model
j + 1), (1813)
where K is the moment of inertia of the molecule. At the same time
the statistical weight of the level e, is g/ = 2j + 1 (provided the
nuclei have no spins).
1 Occasionally, more rigorous expressions are used in which the rotational and
vibrational energies cannot be separated (compare section 119). This is, however,
unnecessary as far as the accuracy of caloric measurements is concerned. In addi
tion to the items of eq. (18.11), sometimes the energy of electronic configurations
(excited states) is also of importance.
300 TEXTBOOK OF THERMODYNAMICS XVIII 117
Two cases must be distinguished with respect to the values of
the quantum integers j. (A) When the molecule is asymmetrical, con
sisting of two different atoms, it can assume all integral values
(j as 0,1,2, . . .). (B) When the molecule is symmetrical, built up of
two identical atoms, only even integral values are permissible 1
(j = 0,2,4, . . .) This distinction was recognized even in the old
quantum theory and was there interpreted as follows: the asym
metrical molecule must be rotated through the angle 2ir until it returns
to the initial position; the cycle of the symmetrical molecule is only
half as large, for after being turned by the angle TT it is in a position
which cannot be told from the initial. The interpretation of quantum
dynamics is somewhat different: when we have two identical atoms
(without nuclear spins) they obey the statistics of EinsteinBose 2 and
admit only quantum states with symmetrical wave functions; these
happen to be those with even numbers j. No such restriction exists
for the asymmetrical molecule.
In either case, we can express the sum of states (18.03) by the
single formula (j = 1,2,3. . .)
(2aj + V exp l Tfl tffo' + *>]' < 18  14 )
r = h 2 /Sir*KkT, (18.15)
if we introduce the symmetry number <r, which is, respectively, equal to
1 and 2, for asymmetrical and symmetrical molecules.
For high temperatures, r is very small and (18.14) yields the equipar
tition value. In fact, putting n = in eq. (18.22), below, we find
in this case
Y r  1/rcr
or log Y r =  log (1/r) + const. Hence (18.07) gives c r = R,
corresponding to the two degrees of freedom of the dumbbell model.
On the other hand, for very low temperatures, only the first two terms
of the sum are significant, log Y r = (2<r + 1) exp [ TO(<T + 1)],
whence c r  R(2a + IJo^cr + 1)V exp [T<T(<T + 1)]. Therefore, the
formula (18.14) represents a gradual rise (with temperature) of the
rotational molal heat from zero to the equipartition value, a behavior
which is in qualitative agreement with the observed facts.
1 We consider here only nuclei without spins. The case of atoms with nuclear
spins will be treated in the next section.
1 Nuclei without spins are, necessarily, built up of an even number of primary
particles.
XVIII 117 THEORY OF SPECIFIC HEATS 301
When quantitative accuracy is desired, the dumbbell model often
turns out to be oversimplified. For one thing, nuclear spins of the
two atoms will introduce changes in the statistical weights, as we
shall see in the next section. For another, the molecule may be
capable of several electronic configurations (excited states) causing a
multiplication of energy levels. The spectroscopy of the band spectra
makes available a material of large and rapidly increasing volume
relating to the rotational and vibrational energy states. For a great
many gases the levels e/ have been accurately measured and tabulated.
If a simple formula is not available, the procedure usually followed is
to substitute into the sum of states (18.03) the spectroscopically
measured levels (or exact theoretical expressions where they are known)
and to obtain the specific heats by numerical calculation.
In view of this, we shall mention here specifically only a particu
larly simple case, that of the doublet structure of the energy levels. It is
not restricted to the rotational energy of molecules but has a quite
general application and was first treated by Schottky. 1 Suppose that
the energy levels occur in pairs ej and ej + Ae (the increment Ae being
constant and independent of /) with the relative statistical weights
go : gi. Every term of Y (18.03) will then appear to be multiplied by
the same factor Y d = go + gi exp ( Ae/jfer). Writing for short
9 = Ae/fc, 7 = \ log (gi/go), (18.16)
we find
Y d = go + gi exp (8/r) = 2go exp (jy  j cosh ^  T j.
Correspondingly, log Y will contain the additional term
log Yd =  + log cosh "" 7 + * log ( 4 S<*i),(18.17)
whence the doublet heat is calculated from (18.07)
(6/2D 2
cosh 2 rz; 
This expression is valid, of course, even when no other energy
spectrum exists, i.e. when each particle is capable only of the two
energy states eo and eo + Ae. In the special case of the rotational
doublets, Ae is not strictly constant. Its formula was worked out by
1 W. Schottky, Phys. Zs. 23, p. 448, 1922.
302 TEXTBOOK OF THERMODYNAMICS XVIII 117
Hill and Van Vleck 1 and shows a slight dependence on j. It seems,
however, that in most cases the simple formula (18.18) gives an
approximation, sufficient within the accuracy of heat measurements
(section 119). The presence of the term c d produces a characteristic
hump in the specific heat curve as exemplified in Fig. 54. The condi
tions are not very different when each energy level, instead of splitting
into two components, splits into several or many closely spaced sub
levels (multiplet structure): the analytical expressions are more
complicated but the result is a hump in the curve of specific heat not
unlike the doublet hump.
It is well known that the energy levels of the linear oscillators are
given by the formula
e,i = hv Q (v + J), v = 0,1,2, . . .
VQ being its characteristic frequency. The vibrational levels of
diatomic molecules are not quite so simple as that but in most cases
they can be sufficiently well represented by an expression of the form
+ i)  *(v + ) 2 L (18.19)
(anharmonic oscillator) where x is a constant. A theoretical founda
tion for this formula was given by Kratzer 2 and in a different way by
Morse. 3 We repeat that, strictly speaking, the vibrational energy is
not entirely independent from the rotational but there exist mixed
rotovibrational terms, which are, however, negligible as far as the
computation of specific heats is concerned. The fact that e/ and e v
are simply additive means that, by taking into account the levels
(18.19), each term of the sum F r (18.14) is multiplied by the same
expression F,<. Since the statistical weight of each vibrational energy
level is 1, we can write
F. 53 exp J   [(v + i)  *(o + *)] J . (18.20)
In other words, the rotational and vibrational sums of states are
multiplicative, Y = 7 r  F v *, whence log Y = log Y T + log Y vi and
c = c r + c vi . The rotational and vibrational heats are additive,
within our approximation, and can be evaluated separately.
It must be noted that the sum of states with the energy levels
* E. L. Hill and J. H. Van Vleck, Phys. Rev. 32, p. 250, 1928. On the basis of
thia work a correction to (18.18) was calculated by G. Gregory (Zs. Physik 78, p. 789,
1932).
A. Kratzer, Zs. Physik 3, p. 289, 1920.
Ph. M. Morse, Phys. Rev. 34, p. 57, 1929.
XVIII 118 THEORY OF SPECIFIC HEATS 303
(18.19) would be divergent if v could increase to infinity. This means
either that the number of states is finite, being cut short by the disso
ciation of the molecule, or that the expression (18.19) represents an
approximation valid only for low order terms. There is no practical
difficulty connected with this because the terms decrease rapidly
even at high temperatures and no ambiguities arise in the numerical
calculations. For a few gases the expression (18.19) does not hold
with sufficient precision; recourse is then taken to the empirical
energy levels tabulated by spectroscopists. On the other hand, when
the coefficient x is small compared with 1 , it is possible to obtain a fair
approximation neglecting it altogether, and extending the summation
to infinity. 1 In this case it can be readily carried out and gives
the result
Y vi = 1/2 sinh (0/2r), 1
f (18.21)
c vi = *0 2 /4r 2 sinh 2 (e/2r), J
where the "characteristic temperature" 9 is an abbreviation for
6 = hvo/k. At any event the law of equipartition of energy does not
apply to expressions of the type (18.19) or (18.20). The vibrational
specific heats, therefore, rise steadily without reaching a limiting
value even at the highest attainable temperatures.
The constant hvo in (18.20) is always considerably larger numeri
cally than rkT of (18.15). Therefore, the vibrational heats come into
play at much higher temperatures than the rotational. The gases
mentioned in the beginning of this section as apparently confirming
the classical theory are those for which the rotational degrees of
freedom have reached complete equipartition at room temperature
while the vibrational degrees are not yet appreciably excited.
Exercise 104. Take a sum slightly more general than (18.14)
fa + n + $) exp [rfo* + n)(*j + n +1)] A/.
yo
A/ = 1 is the increment which j receives from term to term. Write \/r <rj #/;
if T is very small, Ax/ = \fc <rbj is also very small, and the summation can be replaced
by an integration with respect to */. Show that the result of the integration is
Y = (1/ar) exp [ n(n + l)<r], or when a is extremely small,
Y  1/crr. (18.22)
118. Hydrogen (Protium, E^H 1 ) The decline of rotational
specific heats at low temperatures was first observed in hydrogen. 2 Its
1 W. Nernst and K. Wohl, Zs. techn. Physik 10, p. 611, 1929.
* A. Eucken, Preuss. Akad. d, Wise., p. 141, 1912.
304 TEXTBOOK OF THERMODYNAMICS XVIII 118
complete theory was developed in slow steps which were highly
instructive. The simple dumbbell model is here inadequate because
of the nuclear spin of the hydrogen atom. There exist two modifica
tions of the molecule: parahydrogen in which the spins of the two
nuclei are antiparallel, and orthohydrogen in which they are parallel.
The total spin of the parahydrogen is zero and it is indifferent to a
magnetic field, while orthohydrogen is capable of assuming in such a
field three quantized orientations. Therefore, the factors (2j + 1) in
the sum of states (18.14) represent correctly the statistical weights of
those energy levels which belong to parahydrogen while those corre
sponding to orthohydrogen must be multiplied by three. It happens
that the levels of the two modifications divide neatly, parahydrogen
being capable only of even quantum numbers j, and orthohydrogen
only of odd. The corrected sum of states as first given by Hund l is,
therefore,
F= F, + 3F , (18.23)
F,  (4j + 1)<T* 2 ' (2 ' +1) , F a  W + 3)<T' (2 ' +1)(2 > +2) . (18.24)
This is in complete agreement with the band spectrum data where
the lines corresponding to odd levels are three times as strong as those
corresponding to even levels. Hund gave also the theoretical explana
tion of the selectivity of the two modifications with respect to even
and odd levels. It lies in the requirement of wave mechanics that the
wave functions representing the states of the molecule must be anti
symmetric with respect to the two nuclei (compare section 104).
These wave functions consist of two factors which characterize,
respectively, the spins and the rotational motions. It is clear, there
fore, that the arrangement of the spins will influence the selection of
the rotational states. When the spins are antiparallel, the interchange
of the two nuclei produces the same effect as a reversal of the spins:
parahydrogen is, therefore, antisymmetric in the spins and must
have a symmetric rotational function (which happens to be the one
with even j). On the other hand, when the spins are parallel an inter
change of the nuclei does not produce any effect: orthohydrogen is
symmetric with respect to the spins and has an antisymmetric rota
tional function (the one with odd j). We shall, therefore, drop the
prefixes para and ortho and speak instead of symmetrical and anti
symmetrical hydrogen (meaning symmetry with respect to the rota
tions).
F. Hund, Zs. Physik 42, p. 93, 1927.
XVIII 118 THEORY OF SPECIFIC HEATS 305
Yet the specific heats calculated from the sum of states (18.23) do
not agree with those obtained from measurements. 1 The reason for
this has been cleared up by Dennison 2 and lies in the fact that the
two modifications of hydrogen are very slow in reaching equilibrium.
The theory of wave mechanics permits one to foresee that transitions
between the two modifications of hydrogen are extremely unlikely,
and this is borne out by the complete absence of band lines correspond
ing to such transitions. When heat is added to hydrogen, it imme
diately distributes itself over the rotational levels within each modifi
cation separately, but it takes a very long time for the equilibrium ratio
of the two kinds of molecules to be established. The number of mole
cules of symmetrical hydrogen is proportional to F t , that of antisym
metrical to 3F ; we give the equilibrium ratio of these numbers for
several temperatures using the value <rT = 82.6 deg. which follows
from the spectroscopical value 3 K = 4.80 X
Temperature F, : 3Y a
T = oo 1.00:3.00
293. IK 1.00:2.98
78 1.00:1.07
20.4 1.00:0.0145
While at high temperatures symmetrical hydrogen accounts for
only onequarter of the gas, it is practically pure in equilibrium at the
temperature of boiling hydrogen. Under usual conditions, the deter
mination of specific heats takes only a fraction of an hour, and the
actual ratio F, : 3 F in the gas is that corresponding to room tempera
ture and not to the temperature of measurement. In other words,
symmetrical and antisymmetrical hydrogen behave like two inde
pendent gases in the permanent ratio of about 1:3. The appropriate
procedure is, therefore, to calculate separately, by means of eq. (18.07),
the specific heats of the two modifications: c from F,, and c a from F a .
The rotational heat of the actual gas is then
c r  \c. + \c a . (18.25)
The result is indeed in excellent agreement with experiment, as will
1 Experimental determinations were due to: F. A. Giacomini, Phil. Mag. 50,
p. 146, 1925; J. H. Brinkworth, Proc. Roy. Soc. (A) 107, p. 510, 1925; Partington
and Howe, ibidem, 109, p. 286, 1925; Cornish and Eastman, J. Am. Chem. Soc. 50,
p. 627, 1928; Scheel and Reuse, Ann. Physik 40, p. 473, 1913; A. Eucken, Sitzungs
ber. Berlin, p. 141, 1912.
* D. M. Dennison, Proc. Roy. Soc. (A) 115, p. 483, 1927.
Compare: R. T. Birge, I.C.T. V, p. 409, 1929.
306
TEXTBOOK OF THERMODYNAMICS
XVIII 119
be seen from Fig. 49, where the dashed curve is calculated from eq.
(18.25).
An investigation by Bonhoeffer and Harteck l showed that, under
ordinary conditions, the time necessary for reaching equilibrium of
para and orthohydrogen is about ten months. However, it can be
enormously shortened, in fact, reduced to a few minutes, by letting the
gas be absorbed in carbon. In this way it is possible to prepare almost
* 6
100' K 200 300
FIG. 49. Specific heats of hydrogen and deuterium.
pure symmetrical hydrogen which keeps for several months when stored
in a glass container.
119. Deuterium, N, O2> Ck, NO, Cl. Limitations of space do
not permit us to give an exhaustive account of the work on specific
heats of gases, and we restrict ourselves to a few typical examples.
As to rotational heats, the situation in hydrogen repeats itself in
many other gases. Most of them possess a symmetric and an anti
symmetric modification of the molecule, so that the complete sum of
states is
Y r = g.Y. + gaY a . (18.26)
The ratio of statistical weights g : g is, in general, not the same
as in hydrogen (where it is 3 : 1). The two modifications interact but
little and behave, under the conditions of measurement, like two
1 Bonhoeffer and Harteck, Sitzungsber. Berlin, p. 103, 1929; Zs. phys. Chem.
(B) 4, p. 113, 1929.
XVIII 119 THEORY OF SPECIFIC HEATS 307
independent gases. The experimental specific heats are, therefore,
represented by the formula
cr  g ' C t gaC * (18.27)
The vibrational heats are usually calculated from expressions of the
sum of states of the form (18.20).
(1) Deuterium (H 2 H 2 ).
The nucleus of deuterium (deuteron or the heavy isotope of hydro
gen) possesses a double spin (with the quantum number 1, as compared
to ^ of the proton). As shown in the quantum theory, the molecule is
then capable of 6 symmetrical and 3 antisymmetrical spin configura
tions. Moreover, the deuteron obeys the EinsteinBose statistics,
being built up of an even number (2) of protons. Therefore, the
symmetrical spin states combine with the symmetrical rotational
states, etc., and we have g = 6, g a = 3. The specific heats of deuterium
have been calculated with great accuracy (even taking into account
mixed vibrorotational terms) by Johnston and Long. 1 Their results
are represented by the solid curve of Fig. 49. The decline takes place
at lower temperatures than in the case of protium, caused by the,
roughly, two times larger moment of inertia. The dotted curve of
Fig. 49 refers to the mixed protondeuteron molecule rPH 1 as given
by the same authors.
The data given below for several gases are mostly taken from
Trautz and Ader, 2 in whose paper can be found the references to the
older work on the subject.
(2) Nitrogen, N 2 (Fig. SO).
The rotational heat (determined by K = 39.65 X lO" 40 , <rT
= 2.873, g a : g. = 2 : 1) reaches the full equipartition value at 29 abs.
The vibrational levels are represented by 2t 3
0,< = fc 1 **  3374.24(0 + J)  20.66(t> + ) 2 . (18.28)
(3) Oxygen, O 2 (Fig. 51).
1 H. L. Johnston and E. A. Long, J. Chem. Phys. 2, p. 389, 1934.
2 M. Trautz and H. Ader, Zs. Physik 89, pp. 1, 12, 15, 1934. The experimental
data included in Figs. 50, 51, 52 are taken from the following papers: Eucken and
v. Liide, Zs. phys. Chem. 5, p. 413, 1929; Eucken and Miicke, ibidem 18, p. 167,
1932; P. S. Henry, Proc. Roy. Soc. 133, p. 492, 1931; Shillings and Partington,
Phil. Mag. 9, p. 1020, 1930.
1 Compare also: Giauque and Clayton, J. Am. Chem. Soc. 55, p. 4875, 1933.
308
TEXTBOOK OF THERMODYNAMICS
XVIII 119
The equipartition of the rotational heat is reached at temperatures
still lower than in the case of nitrogen. The vibrational levels are 1
2310.7(0
 16.26(o
(18.29)
1.5
1.0
0.5
(
+ (
(
"" x i
calc.) F<
calc.) Fc
obiJEu
obs.) H<
obs.) Sh
I
1
jrmula <
>r mula (
ckn am
nry.
illings ai
'artmgto
18.28)
18.21)
J Miickt
id
i
/
^
^
^^^
/
o
c x
 e vl in Nitrogtn 
200 400 600 800 1000 1200 1400 1600 1800
FIG. 50. Vibrational specific heat of nitrogen.
The specific heat of O 2 has been calculated also by Johnston and
Walker, 2 who used the rigorous spectroscopic expressions for the
energy levels (including the mixed vibrorotational terms). Their
o 15
t"
0.5
^
Z^Z^
**
f
/
S*
o
y
/
+ (
c vl in 2
calc. ) Formula ( 18.29
calc.) Formula (18.21
calc.) Johnston and W
obs.) Euckenand Muc
obs. > Henry
\
)
ilktr
ke
/
/
(
(
(
200 400 600 800 1000 1200 1400 1600 1800
Tabs *
FIG. 51. Vibrational specific heat of oxygen.
data are represented by the dashed curve in Fig. 51. It will be seen
that the difference is appreciable only at very high temperatures.
(4) Air (Fig. 52). 1
The constitution of air is 78.06% N 2 , 21.0% O 2 , 0.94% Ar.
(5) Chlorine, C1 2 (Fig. S3). 1
1 Trautz and Ader, he. rit.
1 Johnston and Walker, J. Am. Chem. Soc. 55, pp. 172, 187, 1933.
XVIII 119
THEORY OF SPECIFIC HEATS
309
The conditions in Cfo are complicated by the existence of the two
isotopes Clasi Cls7 which form three kinds of molecules. The band
spectra, however, show that the amount of ClsyCls? is small and
that chlorine can be taken as consisting of 60% ClasCbs and 40%
1 _
^
10.5
200 400 600 800 1000 1200 1400 1600
Tabs >
FIG. 52. Vibrational specific heat of air.
Theor
(Exp.)
(Exp.
c vj in i
etical
Eucken
Henry
Mr
and Mt
ickt
/
/
^^
]/
/
/
The rotational heat of chlorine is of only academic interest
as it reaches equipartition at 1.5 abs. The vibrational levels of
are given by 2
= 807.81(t;
(18.30)
c v , in CI 2
<cak.) Formula (18.30)
(obs.) Eucken and Hoffmann
+ (calc.) Formula (18.21)
500 1000 1500 2000
FIG. 53. Vibrational specific heat of chlorine.
In the case of ClasCla? the two coefficients must be multiplied
(36/37)* and (36/37), respectively: 2
796.26(0
 5.57(*
(18.31)
1 A. Elliott, Proc. Roy. Soc. (A) 127, p. 638, 1930.
* Trautz and Ader, loc. tit.
310
TEXTBOOK OF THERMODYNAMICS
XVIII 119
(6) Nitric oxide (NO) does not offer much new with respect to the
rotational and vibrational heats. In fact, c r reaches equipartition
below 30 abs. and comes into play only above room temperature.
In the intermediate range NO is interesting because it offers an example
of doublet heat c* as given by the formula (18.18). From spectro
scopic data the doublet difference is Aeo = 354 cal/mol, whence
9 178 deg. As mentioned in section 117, it is not strictly constant.
The measurements were carried out by Eucken and d'Or, 1 who also
calculated the theoretical values from the approximate formula (18.18),
with 7 = 0. Accurate calculations of the specific heat of NO were
carried out by Johnston and Chapman, 2 who used the rigorous
spectroscopic expressions of the energy levels. We give the values of
c d inferred from their results in the last column of Table 41. It will
be seen that the approximation given by formula (18.18) is pretty
good and certainly sufficient in view of the not very accurate experi
mental data.
TABLE 41
SPECIFIC HEATS OF NO
T
cd/R
(obs.)
ca/R (calc.)
Eucken and
d'Or
Johnston and
Chapman
50. OC
0.34
0.325
75.0
....
0.44
0.429
100.0
....
0.39
0.395
127.0
0.27
0.31
0.314
133.3
0.31
0.29
0.295
150.0
0.28
0.25
0.258
173.8
0.22
0.20
0.21
193
0.17
0.17
0.18
218
0.14
0.14
0.15
298.1
0.14
0.09
0.10
(7) Atomic chlorine (Cl) offers another example of doublet heat
(18.10). From spectroscopical data Ae = 2 X 580 cal/mol, 9 = 1299,
7 = 2 log. The calculation 3 leads to the curve shown in Fig. 54.
There are more gases whose complete curves of specific heats
have been accurately calculated, but limitations of space do not
1 A. Eucken and L. d'Or, Nachrichten Gdttingen, p. 107, 1932.
* Johnston and Chapman, J. Am. Chem. Soc. 55, p. 159, 1933.
* W. Nernst and K. Wohl, Zs. techn. Physik 10, p. 608, 1929.
XVIII 120
THEORY OF SPECIFIC HEATS
311
permit us to extend the number of examples, and we can only give
here an incomplete list of references: HgO has been treated by Trautz
and Ader (loc. cit.) ; CO and N2 by Johnston and Davis (J. Am. Chem.
Soc. 56, p. 271, 1934); OH by Johnston and Dawson (ibidem 55,
p. 2744, 1933). The work previous to 1930 can be found in I. C. T. and
in other Tables of Constants.
Figures 5053 contain also the values calculated by Nernst and
Wohl (loc. tit.) with the help of the approximate formula (18.21).
500 1000 1500 2000 2500 3000 3500
T e abs >
FIG. 54. Doublet specific heat of atomic chlorine (theoretical).
The very good agreement shows that this simple way of computation
can be used to great advantage when the coefficient x of the formula
(18.19) is small compared with unity (and, presumably, T not too
high). The values of the characteristic temperature 0, for the common
diatomic gases, as used by them, are given in Table 42.
TABLE 42
CHARACTERISTIC TEMPERATURES
Gas
e
Gas
e
Gas
e
Gas
e
Hi
5950
o a
2220
HF
4130
Li 2
495
Fa
1610
S 2
1030
HBr
3660
Na 2
225
Cl*
888
SC2
570
HO
5110
K 2
130
Br
465
Te 2 . . . .
360
NO
2680
NaK
175
Fa
305
Ni
3374
CO
3060
C1J
545
120. The chemical constant of gases. The notion of the chemical
constant was introduced in section 95. It is that part of the entropy
of a perfect gas which is left undetermined by the first and second laws
312 TEXTBOOK OF THERMODYNAMICS XVIII 120
of thermodynamics. Its determination completes the definition of the
thermodynamic functions, especially of the thermodynamic potential
<p = u Ts + pv t and so makes definite the equations of equilibrium
(6.50). In section 95 we showed that it has the SackurTetrode
expression (14.12) in the case of molecules having only translational
degrees of freedom. Generally, the thermodynamic potential of a non
degenerate perfect gas can be divided into two parts
where <p t refers to the translational and <p* to the inner (rotational,
vibrational, etc.) degrees of freedom of the molecules. To <p t applies the
expression which was derived for a monatomic gas consisting of mate
rial points, i.e. (5.41), (5.43) with c p = SR/2. On the other hand, the
inner motions do not contribute to the term pv and consequently <?, is
identical with the work function ^ whose connection with the sum of
states is given by (18.08):
^ = u Qt + RT[lo% p + f log T  log Y  j t ], (18.32)
where jt is the expression (14.13).
(A) Monatomic gases do not possess rotational or vibrational
degrees of freedom. The sum of states Y = Y e refers here only to the
energy levels of electronic configurations. The term jRriog Y e
contains, however, a contribution to the zero point energy. In fact,
factoring off the first exponential in the sum (18.03) we may write
Y. = exp ( eo/kT) g, or log Y. =  (eo/kT) + log g e . The quantity
gt can be thrown into the form
< exp (GOT, (18.33)
i
where 6f = (e, eo)/. Consequently,
V  uo + RT(lo%p  flog T  j),
jio   1.589 + f logio M + logio g.> (18.34)
(provided the pressure is expressed in atmospheres). o represents
the statistical weight of the fundamental state due to the moment of
momentum of the atom. The quantities Qf are usually so large that
the higher terms of (18.33) are negligible even at fairly high tempera
tures. Only in rare cases are the terms with / = 1,1 2 appreciable.
Data relating to a number of atoms are contained in Table 43. 1
^ Taken from K. Wohl, LandoltBornstein, Second Supplement, p. 1254, 1930.
XVIII 120
THEORY OF SPECIFIC HEATS
TABLE 43
313
Atom
Symbol
of level
*
ef
Atom
Symbol
of level
2
ef
Noble gases
i5
1
'Pa
5 1
H, Cu, Ag, Au,
1 2 5i
2
'Pi
3 (9)
228
alkali metals
J M
'Po
lj
324
Zn, Cd, Hg,
I 'So
I
F
*Pfc
4
earth alkalis
f **
2 Pj4
2
580
C
P
9
Cl
2 P&
4
Si
'Po
1]
2 Pn
2
1260
'Pi
3 (9)
110
Br
2 P%
4
'P a
5J
320
2 P^
2
5270
Sn
'Po
1
J
2 P^
4
'Pi
3
2420
2 P^
2
10900
'Pj
5
4900
Fe
6 I>4
9
Pb
'Po
1
5 #,
7
595
'Pi
3
11200
6 Z?2
5
1010
'P*
5
15200
6 Z?i
3
1270
N,P,As,Sb,Bi
^
4
6Z?0
1
1400
(B) Diatomic gases. The concept of the chemical constant can be
extended to diatomic and polyatomic gases only within the approxi
mation with which log F r , pertaining to the rotational states, can be
regarded as additive to log Y due to the other degrees of freedom.
We have seen in the preceding section that this approximation is,
indeed, quite sufficient for thermodynamical purposes. The range
of temperatures is restricted to those above equipartition of the
rotational energies when Y r = (l/or), according to (18.22) or
Y r = 
+ log K  log <r + log T.
In practice, moreover, it is assumed that the molecule is in its
fundamental electronic state, so that g e reduces to go and the remainder
of log Y becomes log go + log Y vi = log go  t*t/RT. Eq. (18.32),
therefore, takes the form
*  o< + *.< + RT[logp  log T  j],
where ^* can be calculated either from the approximate equations
(18.21), with the data of Table 42, or from the more accurate formula
(18.20). Since the numerical value of logio(h?/&ir*k) is 38.402, we
find for the chemical constant the expression
jio 36.813 +  logic M + logio K  logic * + logic go (18.35)
314
TEXTBOOK OF THERMODYNAMICS
XVIII 120
TABLE 44
CHEMICAL CONSTANTS OF DIATOMIC GASES
Gas
10" K
a
go
j
Gas
lO 40 ^
<T
So
3
H a
0.480
2
1
3.355
HF
1.35
1
1
1.10
F a
25.3
2
1
0.285
HC1
2.646
1
1
0.419
Cla
114
1.57
1
1.45
HBr
3.303
1
1
0.197
Br 8
330 ?
1.42
1
2.49 ?
HJ
4.309
1
1
0.609
Ja
742.6
2
1
2.995
HO
1.500
1
2(0 1 e = 173)
0.561
cij
575 ?
1
1
2.89 ?
CO
14.9
1
1
0.159
2
19.3
2
3
0.531
NO
16.35
1
2(6^ = 182)
0.862
S 2
67 ?
2
3?
1.08 ?
Na 2
179.5
2
1
1.262
Tea
860 ?
2
3
3.09 ?
K 2
184
2
1
1.619
N,
13.8
2
1
0.175
NaK
66
1
1
1.324
The fractional symmetry numbers (for C\2 and Br2 in Table 44)
represent the mean values for the isotopes.
(C) Polyatomic gases. Without entering into the derivation we
only state here the final result for the expression of the thermodynamic
potential *
9 = o< + fc< + RT(\ogp  41og T  j],
jio = 56.563 +  log p +  log 'K  log <r + log go, (18.36)
where ~R = (^1X2^3)^ is the geometrical mean of the three moments
of inertia possessed by the molecule.
A few numerical values are given in Table 45.
TABLE 45
CHEMICAL CONSTANTS OF POLYATOMIC GASES
Gas
j
Gas
j
CO,
0.73*0.05
H*O
1.91
N 2
0.68
NH,
1.66
C 3 H
0.008
CH 4
1.94
In conclusion it must be pointed out that the above expressions
are still incomplete in so far as they do not contain the statistical
weights g n of the nuclear spins. Strictly speaking, one should add to
A. Eucken, Phys. Zs. 30, p. 118, 1929; 31, p. 361, 1930.
XVIII 121 THEORY OF SPECIFIC HEATS 315
the expressions (18.34), (18.35), and (18.36) the term log g n . However,
at temperatures at which the symmetrical and antisymmetrical
modifications of the molecules (sections 118, 119) have practically
reached the ratio g a : g a , corresponding to T = oo , this term is the
same in the molecules and in an equivalent number of free atoms so
that it drops out of the equations of chemical equilibrium. At room
temperature this condition is satisfied for all gases except hydrogen,
for which it is true from about 70 C on. In Chapter XIX we shall
discuss problems of transmutation of matter in which the term log g n
may play a role. We mention, therefore, that in the atom of ordinary
hydrogen (protium) g n = 2, in the deuteron g n = 3, in the helium
atom g n = 0.
121. Heat capacity of a reacting gas mixture. The diagrams of
section 119 give the specific heats of several gases up to temperatures
of 2000 abs. Under these circumstances dissociation may be already
appreciable and must be taken into account. It will, therefore, be
well to discuss how chemical reactions, taking place in a gas mixture,
affect its heat capacity. We shall restrict ourselves to conditions
under which the equation of perfect gases pv = RT gives a sufficient
approximation, permitting us to use the equilibrium theory of Chap
ter VIII.
Let us consider a mixture composed of the gases 1, 2, ... in the
respective mol numbers Ni, N2, . . . N ft , and let us suppose the
following reactions to be possible in the mixture
v\G\ + . . . + vpGp = 0,
0,
etc.
(18.37)
The condition of equilibrium for the first reaction is represented by
the mass law (8.18)
X) "i log xi log K(p,T) (18.38)
i
Substituting the definition of the mol fraction xi = Ni/N (where
N = NI + . . . + Np), we can write
i
i log NivlogN* log K, (18.39)
with the abbreviation v v\ + . . . + v ft . According to eqs. (8.19)
and (8.21)
~, (18.40)
316 TEXTBOOK OF THERMODYNAMICS XVIII 121
Q being the heat of the reaction
I
In the same way we find for the other reactions
2>, log Ni  v' log N = log K'> (18.42)
1 m^
and so on.
We are now ready to calculate the heat capacity at constant pres
sure. The total heat function of the gas mixture is
, fi Xl,
and according to (3.26)
The mol numbers change only inasmuch as the molecules of the
mixture take part in the reactions (18.37). Each of the variations
dNi, therefore, must be expressible in the form
dNi = vi da + v'i da' + . . . , (18.45)
where the quantities a, a', ... are independent of the coefficients v.
Hence
Noting (18.46) and the relation (dxi/d^) p = c p i and writing for
short
2Nic plt (18.47)
11
we have
In the case of a mixture of neutral, nonreacting gases this reduces
simply to
C p =",Nic P i.
XVIII 121 THEORY OF SPECIFIC HEATS 317
The process of change of the mol numbers is supposed to take place
in such a way that the equations of equilibrium (18.39) and (18.42)
are always fulfilled. We can, therefore, obtain (da/dT) p by differ
entiating the first of them and taking into account that the contribu
tion of this reaction to (dNi/QT) p is the first term of (18.46), namely,
.
With the abbreviations
etc., we obtain as the final expression
c p = N [r p + F( X ) ^ + TO ~z +
The mol fractions x refer here, of course, to the state of chemical
equilibrium. They can be obtained either theoretically, by the solution
of the simultaneous eqs. (18.38), (18.42), etc., or empirically from
measurements of the density and other properties of the gas mixture.
For the heat capacity at constant volume, we must start from the
internal energy of the mixture
U = ^,NiU h (18.51)
i
whence
For dNi we have again to use the expression (18.45). Noting that
(dui/dT) v = Cii, 2viUi = AC/, and introducing the abbreviation
i^ (1853)
j
we obtain
In view of the relation (8.19), log K = log K p (T)  vlogp, the
equation of equilibrium (18.39) can be transformed by substituting
for the pressure its expression from the equation of state, p = NRT/V:
Y,n log Ni  log K p  v log T + v log (V/R). (18.54)
2
318 TEXTBOOK OF THERMODYNAMICS XVIII 122
Introducing the abbreviations
/(*) = GE'iV*!)" 1 , /'(*) = (XV' 2 /**)' 1 . ( 18  55 )
and taking into account the relation xi = i + RT, we obtain by
differentiating (18.54) with respect to T
..AJ7
/(*) VaTVV \RT 2 T
The expression for the heat capacity (18.52) is, therefore,
An interesting application of this theory is due to McCollum, 1
who investigated the heat capacity of nitrogen tetroxide (N2O4).
On the one hand, he made calorimetric determinations of C p at temper
tures between 30 and 100 C. On the other hand, he calculated C p
from a formula equivalent to (18.50). The equilibrium in N2O4 was
treated in section 52: the only reaction in it (8.22) has the coefficients
vi 2, V2 = 1 v = 1. As was shown there, the mol fractions can
be expressed in terms of the degree of dissociation as follows:
*i  2{/(l + ), * 2  (1  *)/(! + *) Hence we find for the
(single) functions F(x) and f(x)
D, /(*)  1(1  0/(l + Q(2  Q.
McCollum used empirical values of . The excellent correspon
deiice between measurement and calculation which he obtained is
shown in Fig. 55.
122. Velocity of sound in gases and gas mixtures. Determina
tions of the velocity of sound in gases have been often used as a
simple and, apparently, accurate method of measuring their specific
heats. As was mentioned in section 18, the square of the velocity of
sound in a pure gas is supposed to be given by the formula
> McCollum, J. Am. Chem. Soc. 49, p. 28, 1927.
XVIII 122
THEORY OF SPECIFIC HEATS
319
p being the mass density and M the molecular weight. If the gas fol
lows the equation pv = RT, this expression takes the form (3.40)
a 2 = yRT/n,
1 =
c v
(18.58)
oexp.
theor.
FIG. 55. Specific heat of nitrogen
tetroxide.
It was pointed out that these relations apply without regard to
whether c p , c vt y are constants or functions of temperature.
However, the specific heats c v
measured by the sonic method in
oxygen and in air 1 at high temper
atures (from 100 to 1000 C) are
considerably below those determined
by other methods. In a similar way,
the sound velocities in NO give the
same specific heat at low tempera
tures as at C, 2 instead of the
increase tabulated in Table 41 of
section 119. This fact seemed to
indicate that the changes of state,
produced by the sonic waves, are
too rapid to allow the equilibrium of
the molecular degrees of freedom to
be established. In fact, the disper
sion of sound waves in gases (CC>2) had already been observed a few
years earlier 3 and was interpreted as a lag in the molecular adjust
ment. 4 Extensive work on the subject was done by Kneser, 6 who was
able to show that it is the vibrational energy which requires periods up
to 10~~ 2 sec to assume its equilibrium value, while the rotational
degrees of freedom respond much more quickly. At low frequen
cies the ratio 7 of the formula (18.58) was experimentally equal
to 1 + R/(c t + c r + c*i), at high frequencies, to 1 + R/(ct + c r ).
These investigations were greatly influenced by a paper of Ein
stein's 6 which suggested sonic measurements as a method of deter
1 W. C. Shillings and J. R. Partington, Phil. Mag. 9, p. 1020, 1930.
1 Partington and Shillings, Phil. Mag. 45, p. 416, 1923.
1 G. W. Pierce, Proc. Am. Acad. 60, p. 271, 1925.
4 K. F. Herzfeld and F. O. Rice, Phys. Rev. 31, p. 691, 1928.
*H. O. Kneser, Ann. Physik 11, pp. 761, 778, 1931; 16, p. 337, 1933; H. O.
Kneser and V. O. Knudsen, Ann. Physik 21, p. 682, 1935.
6 A. Einstein, Sitzungsber. Berlin, p. 380, 1920.
320 TEXTBOOK OF THERMODYNAMICS XVIII 122
mining the rates of reaction in gas mixtures. While this paper was
indirectly responsible for the research on dispersion in pure gases,
the results of this research defeated its original purpose. The obser
vations in mixtures are too difficult to interpret when possible abnor
malities of the components themselves must be taken into account. 1
For this reason, we shall calculate here the velocity of sound in mix
tures only for low frequencies when the question of lag does not
arise.
The velocity of sound in a gas mixture of the total mass M can
be written
a 2 = (dp/dp) a = ~ V 2 (dp/dV) s /M, (18.59)
since the mass density is p = M/V. Let us envisage a mixture of a
given initial constitution. If its reacting components are always in
equilibrium, the total volume is completely determined by the tempera
ture and pressure. The system is, therefore, simple in the sense of
section 2: its entropy differential, dS = (dU + pdV)/T, can be
regarded as a function of pressure and volume, and the equation of
the adiabatic can be written
since OX/9F) P = (dU/dV) p + p, in view of the definition of the heat
function X = U + pV. Because of (3.12) and (3.26)
faX
V 1 = *"V ITI / \
\dp/v \dp/v
whence
and
(dp/dV)a BB y(dp/dT) v /(dV/dT) p = 7(9/9tOrl
 (18.60)
7 = Cp/Cv ]
This equation holds generally for any simple system. When we
apply it to a gas mixture, the heat capacities in the ratio y are, of course,
1 C. E. Teeter, J. Chem. Physics 1, p. 251, 1933. This paper contains a good
review of the experimental and theoretical work about measuring rates of dissocia
tion in nitrogen tetroxide by the sonic method.
XVIII 123 THEORY OF SPECIFIC HEATS 321
those calculated in the preceding section. From the equation of state,
pV = NRT,
We transform this with the help of the relations (18.43), (18.46),
and (18.54), obtaining
1 + tf ) + //'(*)
(18 ' 62)
It will be noted that the last factor is much less affected by the
heats of reaction than 7: it is linear in Q/RT, etc., while Cv, C p
contain the squares of these large numbers.
Experimental work concerning the velocity of sound has been
published by several authors. 1 While the results are not quite con
sistent in their bearing on the rate of the reaction and the speed
with which equilibrium is established, it seems that indications of a
lag in reacting gas mixtures are to^be found only in the measurements
of absorption of sound. There is no convincing proof of the existence
of dispersion, and the formula (18.62) appears to represent the observed
velocities fairly well. This would mean that chemical equilibrium is,
practically, reached after the lapse of only 10~ 4 sec.
123. Solids. Einstein's model of a solid body was described in
section 93. It treats the atoms of the solid as harmonic oscillators,
all independent and having the same frequency v. For the energy
levels of the threedimensional harmonic oscillator the quantum theory
gives
e ni n 2 n, = kv(m + W2 + W3 + f),
where n\ t n%, #3 are the quantum numbers of the three degrees of
freedom. The sum of states is
exp [*(i +H2 + H3+ f)] = C (18.63)
\ l ~ e )
(x = hv/k). The energy and entropy can be derived from it by eqs.
1 Compare the review in the paper of Teeter's (he. cto.)
322 TEXTBOOK OF THERMODYNAMICS XVIII 123
(18.06) and (18.09), and are those given in section 93. In particular,
the mean energy of the oscillator is, per degree of freedom,
 "*
' e  l ' 2
As was mentioned before, the agreement of Einstein's formulas
with the measured specific heats is not very close. The reason for this
discrepancy was pointed out in a later paper 1 of his: the model does
not represent the actual conditions accurately enough because, in
reality, the atoms of a crystal do not oscillate independently of one
another with the same frequency but form a coupled system. The
forces of interaction influence the characteristic frequencies so that
they all become different and are drawn out into a spectrum. More
accurate theories were given from this point of view by Debye 2 and
independently by Born and Von Karman. 3 Debye's theory is par
ticularly simple: he assumes that the 3Z frequencies of a body con
taining Z atoms can be calculated as if it were a homogeneous elastic
continuum. In other words, these frequencies correspond to the
fundamental tone and to the first 3Z 1 overtones of its elastic
spectrum. The following considerations support this assumption.
In the region of high temperatures prevails equipartition and the
energy depends only on the number of frequencies (or degrees of
freedom) and not on their values. On the other hand, at low tempera
tures, the high order overtones are not appreciably excited ; according
to eq. (18.64) their mean energy becomes very small. All depends
then on the low order frequencies, and for these the approximation by
a continuum is quite adequate because their wave length is large
compared with the atomic distances. Therefore, we should expect
accurate results both at low and at high temperatures.
The calculations are easily carried through in the case of isotropic
elementary materials. If we take the solid in the shape of a cube with
the edge /, 4 the theory of elasticity gives the following expressions for
the frequencies of the standing longitudinal and transverse waves
v  ai(m 2 + 2 2 + w 3 2 ) H /2/, v = a 2 (ni 2 + n 2 2 + w 3 2 ) H /2/, (18.65)
where a\ and a* are the velocities of the longitudinal and transverse
wave systems. Every triple of integral numbers wi, w 2 , m determines
a state of the body (standing wave) in each of the two wave systems,
1 A. Einstein, Ann. Physik 35, p. 679, 1911.
P. Dcbye, Ann. Physik 39, p. 789, 1912.
*M. Born and Th, Von Karman, Phys. Zs. 13, p. 297, 1912; 14, pp. 15, 65, 1913.
4 R. Ortvay, Ann. Physik 42, p. 745, 1913.
XVIII 123 THEORY OF SPECIFIC HEATS 323
the corresponding frequencies being given by eqs. (18.65). It is easy
to see that the number of all longitudinal states with frequencies
smaller than a given value v is
~
(18.66)
.
where V = P is the volume of the cube. In the transverse system the
number must be doubled: every transverse wave has two degrees of
freedom because of the two possible states of polarization
(18 ' 67)
Exercise 105. If we imagine the numbers ni, n 3 , n 9 plotted in three orthogonal
directions, their integral values will represent a cubic lattice. The number Zi(v) is the
number of lattice points within the surface given by the first eq. (18.65). Show
that the expression (18.66) holds when / is very large. i, n^ n\ are positive.
If the solid were really a continuum, there would be no upper limit
to its characteristic frequencies. But it has, actually, atomic structure,
and the total number of its possible states must be equal to the number
of its degrees of freedom, namely 3Z. Debye, therefore, makes the
assumption that the elastic spectrum is cut off at an upper limit VQ
determined by the condition
3Z.
Substituting (18.66) and (18.67) we find
The number of characteristic states in the frequency interval from
v to v + dv is
QZ
dZi(v) + dZ*(v)  Mv.
v<f
Each of these states is a harmonic oscillation whose mean energy
is given in the quantum theory by the expression (18.64). The con
tribution of the frequency interval dv to the internal energy is, there
fore, QZpZiPdv/vo 3 * The total internal energy is obtained by integrat
ing this expression from to VQ and is, per mol,
" *dv , /40 ^
 + " (18 ' 68)
324
TEXTBOOK OF THERMODYNAMICS
XVIII 123
The specific heat results from differentiation with respect to T
(contained in x). If we introduce the abbreviation hvo/k = 9,
c =
(18.69)
This formula shares with Einstein's (18.64) the property that
the individual characteristics of the substance are all contained in
the single constant 9, the characteristic temperature. If we plot c
against T/9 (the reduced temperature) we obtain for all bodies one
and the same universal curve. In the limiting case of very low tem
peratures (r/9 1) the integral can be regarded as constant and
equal to 47r 4 /15. The formula becomes then
127T 4
c =
The agreement with experiments can be judged from Fig. 56 l
giving the measured c v . The value of 9 for each substance is here
2 3 4 5 6 7 8 9 2 1
Pb, Ag, KCI, Zn, NaCI, Cu, Al, CaF 2 C
FIG. 56. Specific heat of solids.
chosen so as to give the best fit for that substance with the continuous
curve representing eq. (18.69). Curve II is in the normal position;
the others are displaced because the experimental material was too
large to be accommodated in a single curve. The key to the sub
stances is contained in Table 46.
1 Taken from E. Schrodinger, Phys. Zs. 20, pp. 420, 450, 474, 497, 523, 1919.
XVIII 123
THEORY OF SPECIFIC HEATS
325
TABLE 46
KEY TO FIG. 56
Substance
Chemical
symbol
Temperature
range
e
Points in Fig.
I
II
III
Lead
Pb
Tl
Hg
I
Cd
Na
KBr
Ag
Pt
Ca
KC1
Zn
NaCl
Cu
Al
Fe
CaF 2
FeS 2
C
14573
23301
31232
22298
50380
50240*
79417
35873
88
96
97
106
129160
159
177
215
225
230
217.6
235
287
315
392
454
499
645
18602230
X
O
V
D
A
+
V
X
V
D
A
+
V
D
X
+
A
Thallium
Mercury
Iodine
Cadmium
Sodium
Potassium bromide. . .
Silver
Platinum
Calcium
2262
23550
33673
25664
14773*
19773
3295 *
17328
2257 *
301169
Sylvine
Zinc
Rocksalt
Coooer
Aluminum
Iron
Fluorspar
Iron pyrites . .
Diamond
* Rises above the curve after these temperatures.
The formula (18.69) was derived on the assumption that the sub
stance is isotropic and elementary. Within this scope, it represents
the facts on the whole very well although a few exceptions are known.
But Fig. 56 shows that it applies, in some measure, also to substances
of a more complicated structure, presumably because they do not
deviate greatly from isotropy.
The approach of Born and Von Karman permits to treat the more
general cases, at least, by approximation. If the substance is an
element in an anisotropic system of crystallization, the elastic waves
propagate in it with three different velocities and form three sets of
overtones. Born l finds that the specific heat is then the arithmetical
mean of three terms of the Debye type (18.69), each with a different
characteristic temperature. If the substance is a chemical compound
containing r different kinds of atoms, there still exist, in addition to
the spectrum of overtones, r 1 characteristic frequencies represent
1 M. Born, Dynamik der Krystallgitter, 1915.
326
TEXTBOOK OF THERMODYNAMICS
XVIII 123
ing the oscillations of these atoms with respect to one another. The
molal heat is, therefore, expressed by three Debye terms and r 1
terms of the Einstein type (13.25). Unfortunately, these expressions
are so unwieldy that their comparison with experimental results is
difficult.
The third power law at very low temperatures remains valid also
in the general case because the Debye terms satisfy it and the Einstein
terms become negligible. While a great many substances fulfill the
third power law down to the very lowest attainable temperatures, 1
other metallic and nonmetallic solids show marked deviations in the
region of the temperature of
liquid hydrogen and below.
The measured specific heats lie,
in this case, above those pre
dicted by Debye's formula. It
was suggested that in metallic
substances the discrepancy
may be due to the specific
heats of the free conduction
electrons (compare sections 102
and 111). In fact, in the case
of silver the influence of the
free electrons seems well
authenticated by recent
measurements in the laboratory
of Leiden. 2 In Fig. 57 is plot
ted the difference between the
.0008
fWtf%
.uuuo
V0004
/
/^
8
\
\
OrtHO
/
\
.UOU&
c
/
\
t
/
\
T 2 4 6K
FIG. 57. Electronic specific heat in silver.
measured c and the value
predicted by the third power law (18.70). It shows both the linear
dependence required by the Sommerfeld formula for free electrons
(17.34) and the correct absolute value on the assumption that there
is one electron to every atom of silver.
The only other (nonsupraconductive) metal investigated in the
same temperature region is zinc, and here the free electrons do not
give a sufficient explanation of the trend of specific heats. Even
after correcting for them, there remains a humplike residual curve
with a maximum at 4 abs., not unlike in its shape to that of Fig. 54.
Indeed, it was suggested that they are due to the same cause, doubling
of the lowest quantum level of the zinc atom. However, the relation
1 Compare in this connection, the work of Simon and collaborators (Ann. Physik
68, p. 241, 1922; Zs. phys. Chemie 123, p. 383, 1926; Zs. Physik 38, p. 227, 1926).
W. H. Keesom and J. A. Kok, Physica 1, p. 770, 1934.
XVI II 123
THEORY OF SPECIFIC HEATS
327
2.0
1.5
0.5
i
0.4
1.6 2.0
of the position of the hump and its height is such that only a small
fraction of the zinc atoms could possess the double level, a circumstance
which makes the explanation doubtful. There are some solids (such
as gadolinium sulfate, 1 samarium sulfate, 2 orthohydrogen 3 ), in which
the doubling or tripling of the atomic and molecular levels gives a
complete explanation of the anomalies of specific heats according to
(18.18) or to similar formulas. We reproduce here the very instructive
curve measured by Giauque and McDougall in gadolinium sulfate
(Fig. 58).
In other cases (e.g. tin and silicon 4 ) there were observed humps
in the curves of specific heats lying at somewhat higher temperatures
whose detailed explanation is still
lacking. Simon suggested that they
may be due to a doubling of the
energy levels according to the
formula (18.18). It should be men
tioned, however, that electrons
which are not completely free and
cannot cause electric conduction
may nevertheless contribute to the
specific heats, sometimes even in
a larger measure than the free
electrons. At extremely low tem
peratures the specific heat due to
them is proportional to r, in nonmagnetic materials, and to T*, in
ferromagnetic ones. 6 At somewhat higher temperature this cause,
being essentially a multiplet action (compare section 117), should
produce a hump similar to that of formula (18.18).
Recent work by Jaeger and coworkers (on metals at high tempera
tures) drives it home that our theoretical knowledge of specific heats is
still incomplete. Measuring very accurately, 6 they found that c p is
structure sensitive and varies appreciably with the mechanical treatment
of the specimens. Another series of investigations 7 was devoted to
testing Neumann's law (section 16) with the result that it was found
satisfied, within the experimental accuracy of 1%, in mixed crystals
1 Giauque and McDougall, Phys. Rev. 44, p. 235, 1933.
2 Ahlberg and Freed, Phys. Rev. 39, p. 540, 1932.
1 Mendelssohn, Ruhemann, and Simon, Zs. phys. Chem. (B) 15, p. 121, 1931.
4 F. Simon, Sitzungsber. Berlin, p. 477, 1926.
F. Bloch, Zs. Physik 52, p. 555, 1928; Leipziger Vortrage, p. 67, 1930; P. &
Epstein, Phys. Rev. 41, p. 91, 1932.
6 Jaeger, Rosenbohm, and Bottema, Proc. Amsterdam 35, pp. 763, 772, 1932*
7 Bottema and Jaeger, ibidem, 35, pp. 352, 916, 929, 1932.
0.8 1.2
abs
FIG. 58. Anomaly of specific heat in
gadolinium sulfate.
328
TEXTBOOK OF THERMODYNAMICS
XVIII 124
(AgAu). On the other hand, in homopolar intermetallic compounds
(AuSn, PtSn) there were found appreciable deviations from Neumann's
law which, moreover, increased with rising temperature (between
and 200 C).
124. Liquids. The liquid state of matter is the most complicated
and has not yet been made amenable to theoretical treatment. There
exists a number of elements and compounds having in the liquid state
approximately the same molal heats which are characteristic of solids,
according to the laws of Dulong and Petit and of Neumann (section
16), namely, about 6 cal/mol per atom (Table 47).
TABLE 47
SPECIFIC HEATS OF LIQUIDS
Liquid
c
n
c/n
Liquid
c
n
c/n
Hg
6.7
1
6.7
H 2 O
18
3
6.0
N,
13.2
2
6.6
S 2 C1 2
22.4
4
5.6
2
12.8
2
6.4
SiCl 4
34.4
5
6.9
AgBr
14.0
2
7.0
TiCU
36.8
5
7.4
T1C1
13.6
2
6.8
20
19
18
17
16
15
C(p)
in cal /mol
This would indicate that the conditions are similar to those in
solids in that one must take into
account both the kinetic and the
potential energy of the atoms.
However, many liquids have atom
ic heats much larger than 6 cal.
Because of its practical im
portance, much attention was
devoted to the specific heat of
water. Of particular interest is
the dependence of c p on pressure
represented graphically in Fig.
59. These curves were not de
termined by direct measurement
but inferred from the thermal
expansion at different pressures 1
by means of the relation p (4.59),
or Cp = /r(aV3^ 2 )p dp. We
shall see that these data have an important bearing on the question
of specific heats of aqueous solutions.
*Bridgman, Proc. Am, Acad. 48, p. 309, 1913; F. Zwicky, Phys. Zs. 27, p. 271,1926,
FIG.
123456789
p10T 3 inkgXcm 2
59. Specific heat of water as a
function of pressure.
XVIII 124
THEORY OF SPECIFIC HEATS
329
As was shown by Zwicky, the specific heats of nonelectrolytic
solvents in water are additive: i.e. the total heat capacity of the
solution can be expressed by the formula
No c p o + Ni c p i
(18.71)
where C P Q = 18.01 is the molal heat of water (at 15 C) and c p i of the
solute. It is interesting to note that the values of c p \ observed in the
solution are pretty close to those observed in the solid state of the
solute, as appears from Table 48.
TABLE 48
SPECIFIC HEATS OF SOLUTES
Substance
Dextrose
C 6 H 12 6
Lactose
CiaH^Oii
Urea
OCN 2 H 4
Glycerin
C,H 8 0,
Tartaric acid
C 4 H 6 6
c p i from eq. (18.71) ..
Cpi in solid state
55.2
51 7
104.5
98 5
24.8
19 3
56.0
53
67
62
When we turn to aqueous solutions of electrolytes, we meet with
the unexpected fact that the heat capacity is not increased but lowered
by the solutes. A qualitative explanation of this behavior was given by
Tammann, 1 who pointed out that the effect is analogous to a rise of
pressure to about 500 atm: the electrolyte must increase, in some
way, the pressure under which the particles of water stand. Accurate
modern measurements in solutions of alkali chlorides 2 showed that
at low molalities (up to m = 0.5) the heat capacity is still well repre
sented by the linear formula (18.71). However, the coefficients c p \
here take negative values of the order 40.
Taking up Tammann's suggestion, Zwicky (loc. tit.) developed it
into a detailed theory. The field of the ions exercises large ponderomo
tive forces on the electric dipoles of the water molecules which actually
produce a strong compression of the aqueous medium in their vicinity.
Each ion is, therefore, surrounded by a little sphere within which the
specific heat of the water is reduced by pressure. As long as the con
centration is low, the spheres do not overlap and the reduction of the
heat capacity is proportional to their number, i.e. to the molality.
Hence the validity of the linear expression (18.71). The quantitative
calculation of this effect gave the right order of magnitude.
1 G. Tammann, Innere Krafte und Eigenschaften der Losungen, 1907.
1 Richards and Rowe, J. Am. Chem. Soc. 43, p. 770, 1921; 44, p. 684, 1922.
CHAPTER XIX
EQUILIBRIUM INVOLVING RADIATION
125. Equivalence of mass and energy. We devoted Chapters
XIII to XV to questions pertaining to the determination of the abso
lute value of the entropy S. It should not be forgotten, however, that
there is another fundamental function, the internal energy U, whose
absolute value is also indeterminate. As far as the first and second
laws are concerned, S and U are on the same footing since only the
differentials of these two functions are defined in thermodynamics
(compare sections 13 and 22). Both contain, therefore, additive inte
gration constants. That the indeterminacy of the energy constant
was much less emphasized was due to historical and practical reasons.
It was taken as a matter of course, because the thermochemical
methods of measuring the differences of these constants in a reaction
had already been developed before thermodynamics was established
as a separate science.
It became possible to determine the absolute value of the energy
after Einstein l had established the law of proportionality of mass and
energy as an important result of his theory of relativity. The fact
that the inertial mass of a particle depends on its kinetic energy was
not an exclusive feature of relativity but was proved long before its
advent. But only the theory of relativity gave a sharp definition of
the concepts here involved, and abolished the difference between
inertial and gravitational masses. Let M be the mass of a system,
E its total energy, and c the velocity of light. Einstein's law is then
expressed by the simple relation
E Me 2 . (19.01)
If we consider 1 mol of a substance at the temperature T = 0, its
internal energy will be UQ, while its mass will be equal to its molecular
(or atomic) weight ju Einstein's law gives, therefore,
*o => 2 . (19.02)
1 A. Einstein, Ann. Physik 18, p. 639, 1905.
XIX 126 EQUILIBRIUM INVOLVING RADIATION 331
Strictly speaking, the mass depends on the temperature since the
energy changes with it. To be exact we should have specified that M
is the molecular weight at T = 0. Chemistry proper deals with dif
ferences Ao (for instance in dissociation of molecules) so small that
the attendant change of mass AM = Awo/c 2 cannot be detected by the
finest measurements. In this chapter, however, we shall treat proc
esses in which an appreciable part of the energy of the reacting sub
stances is converted into radiation. To the same extent their mass
will undergo a loss: conservation of energy and conservation of mass
hold only for matter and radiation together and not for the material
part of the system separately. Compared with the huge changes of
mass involved in these processes the difference between /u at T =
and T = 300 abs is irrelevant and we shall understand under the
symbol AC the molecular weight as tabulated by international agreement.
As an example we write out the explicit expression for the thermo
dynamic potential of a monatomic perfect gas
V = RT(log p   log T  j) + nc*. (19.03)
For some applications it will be more convenient to express the
pressure in terms of z (the number of molecules or atoms per unit
volume) by the relation (1.15), p = zkT,
V = RT(log z   log T  j + log *) + c? (19.04)
126. Energy and entropy of black radiation. The theory of radi
ation is so extensive a subject that it is best treated in separate books, 1
and it is not advisable to discuss it at length in a textbook devoted to
thermodynamics. We limit ourselves to a brief review of some of its
concepts and laws which shall be needed in the remaining part of this
chapter.
Let us consider a closed system part of which is free of matter
(vacuum) and filled only with radiation, while matter is present in
other parts. Left to itself such a system reaches a state of equilibrium
in which the temperature T is uniform throughout and the radiation
in the evacuated part has quite definite properties. It is called black
radiation, and its spectral composition is given by a universal law,
independent of the nature of the material part of the system in equilib
rium with it: the energy density u/lv (i.e. energy per unit volume)
belonging to the interval of frequencies between v and v + dv is a
universal function of v and T only
M"  F(v,T)dv. (19.05)
1 For instance, M. Planck, The Theory of Heat Radiation, 1914,
332 TEXTBOOK OF THERMODYNAMICS XIX 126
This implies, of course, that the total energy density
/CO /*00
u&  / F(v,T)dv = u(T) (19.06)
^o
is a function of the temperature alone.
If black radiation is enclosed in an evacuated container, it exerts
a pressure p on its walls and through it resists compression. Assuming
the walls to be movable, the radiation does the work p dV when the
volume is increased by dV (compare section 7). If an element of heat
dQ is imparted to the radiation, the first law of thermodynamics
requires that it be converted, in part, into an increase of its energy U,
in part, into work done by it
dQ = dU + pdV.
The second law can be also applied to radiation and includes the
familiar expression (4.07) for the entropy differential
<ZS= (dU + pdV)/T. (19.07)
In particular, when the volume of the system remains constant
(V = const, dV = 0) the increment of the entropy is
dS  dU/T. (19.08)
The same results hold, of course, if the vessel is not quite evacuated
but contains, in addition to the radiation, gases in equilibrium with it,
of so low a density that their refraction is negligible, so that the prop
agation of light is appreciably the same as in vacuo. This is the
case we shall consider in the following sections. The restriction to
low densities, though not absolutely necessary, greatly simplifies mat
ters and saves lengthy discussions.
The general relation (19.08) which contains practically no restrict
ing assumptions about the function U is all we shall need. As far as
our applications are concerned, the rest of this section has a purely
academic interest. The quantum theory of radiation leads, in its
present form, to the result that empty space contains a certain con
stant energy density o even at T = 0. It is usually thought that
this is merely a formal defect and that this zero point energy (which is,
by the way, infinite) has no physical reality. There were, however,
those who argued that a part of it may have physical significance,
that it may influence the absolute entropy constant of the radiation
and through it the equilibrium with matter. We shall see in the
^t section that the conditions of equilibrium are independent of the
XIX 126 EQUILIBRIUM INVOLVING RADIATION 333
absolute entropy of radiation. Nevertheless the question how it is
affected by UQ is worth looking into. We shall write, therefore, for
the energy density, u + Wo, meaning by u that part which is due to
temperature radiation. The total radiation is then U = (u + u$)V\
as to pressure, it is known to be p = w/3, since it is certain from
empirical and theoretical considerations that the zero point energy
does not contribute to it. The entropy differential (19.08) takes the
form
V du /4 \ d V
dS  dT + (j u + no ) (19.09)
The reciprocity relation (2.10) is easily integrated and gives
u = ar 4  f wo, (19.10)
a being a constant. We see that the inclusion of a constant term UQ
(not exerting pressure) into the energy density leads to a contradiction.
The result gives for p = \u \aT* Jwo, so that UQ does, after all,
produce a negative pressure. It follows that the zero point energy,
whether it is in other ways real or not, must be left out in computing
the entropy. We arrive thus at the usual form of the StefanBoltz
mann law
u = ar 4 , (19.11)
and substituting this into (19.09)
5 = ar 3 F+5 .
So is an integration constant and as such must be independent of V
and T. It is obvious that it cannot be anything but So = 0, other
wise we would have still a finite entropy when the volume is reduced to
zero, i.e. when we have no system. The entropy density is, therefore,
* =  = i r3  (19tl2)
Comparing this with (19.11), we verify
ds  du/T. (19.13)
This equation is compatible with the following interpretation form
ing an important part of the theory of black radiation. The entropy
density 5 is considered as the sum of the densities sjiv belonging to
the different frequency intervals
/'
/o
334 TEXTBOOK OF THERMODYNAMICS XIX 127
where $ is defined by
ds v  du,/T. (19.14)
In fact, the integration of this relation with respect to dv (at con
stant T) gives eq. (19.13). The relation shows that the frequency
intervals of the radiation can be regarded as statistically independent
systems in equilibrium with one another at the temperature T (com
pare section 22). In the particular case of constant volume (dV = 0),
we can write
(1915)
ds T 9 (19 * 15)
if we denote U v Vu,, S, = Vs .
127. General conditions of equilibrium. Let us envisage a system
enclosed by rigid, perfectly reflecting, adiabatic walls and consisting
of a sufficiently thin mixture of perfect gases (compare preceding sec
tion) and radiation. Let N\, N%, . . . N ft be the mol numbers of the
gases; the total energy and total entropy of the system are then
if UR, SR stand for the energy and entropy of the radiation. The total
volume of the system is expressed in terms of the molal volume vi of
each gas as
The conditions of equilibrium (compare section 31) are given by S
having its maximum while U and V stay constant, in view of our
assumptions about the walls. In variational form they are
or
JL*
+ 8U* = 0, (19.16)
+ 8S*  0, (19.17)
0, (/ = 1, 2, . . . 0) (19.18)
We can regard the first equation as the main condition, the others
as subsidiary ones. The difference between this system and that of
section 40 lies only in the presence of radiational terms. Apart from
XIX nl EQUILIBRIUM INVOLVING RADIATION 33$
this, the procedure is exactly the same as in treating the ordinary gas
mixture. We first eliminate the variations of the internal energies by
means of (4.15) and (19.13)
dU R = T R dS R ,
whence (19.16) takes the form
+ uitNi) + T R dS R 0. (19.19)
Now we apply the method of Lagrangean multipliers (section 38) :
we multiply the first subsidiary condition (19.17) by Xo and the others
(19.18) by X/, (/ SB 1, 2, . . . 0), and add them to the main condition
(19.19)
#i( Pi
X, vi)*Ni] + (T R + Xo)SS*  0. (19.20)
As was explained in section 38, the variations foj, dS R , dvi can now
be considered as independent, whence their coefficients must vanish
 Xo  Ti = TR  T
(i.e. the temperature is uniform) and
Xz  pi.
The last equation supplies only the physical interpretation of the
multipliers Xj. Recalling the definition (5.33) of the molal thermo
dynamic potential ?i = u\ Tsi + pi vi, 1 we can write what remains
of (19.20) in the form
T] PI#I  0,
which is identical with (6.41) derived without considering radiation.
If the reaction taking place in the mixture is expressed by the equation
0, (19.21)
the condition of equilibrium is
T] n Vi  0. (19.22)
1 As in (8.08) we use the notation "Qi (and not <PI) to indicate that this function is
calculated as if the other gases were not present, i.e. in terms of the partial pressure.
336 TEXTBOOK OF THERMODYNAMICS XIX 128
Though formally identical with (6.49), these conditions have a
wider scope of applications. In deriving them we explicitly admitted
that the energy and mass of the material part of the system need not
remain constant but may be converted, wholly or in part, into radia
tion. We are, therefore, free to apply them also to such reactions
(19.21) in which substances disappear altogether under emission of
radiation (annihilation of matter) or in which one kind of atoms dis
appears and another is created (permutation of elements). We may
also mention that, because of eq. (19.15), we would have obtained the
same result writing in (19.16), (19.17), dU v , 5S,, instead of dU Rt &S Rt
where v is the frequency of the photons emitted in the reaction. In
other words, the reacting substances set themselves in direct equilib
rium with the frequency interval corresponding to their emission and
through it, indirectly, with the rest of the radiation.
Using the expression (19.04) for the thermodynamic potential of
monatomic gases, we obtain from (19.22)
"i log zi = log / + I v log r  Ac 2 /RT, (19.23)
i
where v = v\ + v^ + . . + v$\ the change of mass in the reaction is
and
log / = Z"tf'i  v log k. (19.24)
i
According to section 120, the chemical constant of a monatomic
gas is jio = 4.417 + logio M + log g, where g must include also the
nuclear statistical weight, and logio k = 15.8630. Therefore,
! log Zi = log K n ,
T  ' '   * (19 2S)
logio K. = 20.280? +vi (f logio M< + logio gi)
4.695 X 10 12 A M
+ f v logio T 
128. Negative and positive electrons. Equilibrium of matter and
radiation as a problem of thermodynamics was first treated by
O. Stern. 1 He envisaged a state of affairs in which atoms disappear,
being converted into radiant energy, and others suddenly appear in
the field of radiation, absorbing the amount of energy equivalent to
1 0. Stern, Zs. Elektrochemie 31, p. 448, 1925.
XIX 128 EQUILIBRIUM INVOLVING RADIATION 337
their mass. As science progressed processes of this sort were regarded
as less and less speculative. With the discovery of the positive elec
tron (positron) by C. D. Anderson 1 the creation and annihilation of
electron pairs became an established fact. The tracks of the two
electrons (one positive, one negative) produced by a cosmic ray or
7ray can be observed in the Wilson chamber. There is also direct
evidence that they recombine, emitting their whole energy in the form
of two photons.
If we characterize the electron and positron by the subscripts e and
p, eq. (19.21) of the reaction becomes
G e + G p  0, (19.26)
whence ? = v p = 1, v = 2, and 2
K. (19.27)
For the electron of either sign f log /z + log g = 4.59 (compare
section 113), A/* = 2/x = 1.096 X 10"*
5 146 X 10 9
logio K = 31.38 + 3 logic T    jr  (19.28)
In the empty space, in the absence of all other types of matter,
positive and negative electrons would exist in equal numbers, z p = z et so
that the density of positive electrons would be
z p = K*. (19.29)
This number is exceedingly small at ordinary temperatures and
becomes appreciable only when the temperature approaches 10 8
degrees. In fact, for T = 8 X 10 7 deg, we find z p = 2.5 X 10~ 6 cm" 3 ,
or 1 positron per 40000 cm 3 . From then on the rise is rapid: for
T = 2 X 10 8 , we already have z p = 1.9 X 10 15 cm 3 .
Of course, temperatures of this order do not prevail in interstellar
space. The only place where they may possibly exist is the interior
of stars. Here the conditions are different from those in empty space,
in that there is a large excess of free negative electrons, which depress
the density of the positrons in the way explained in section 53. Let
us denote by ZQ the density of the free electrons which are in the star
from the start, being stripped by ionization off the atoms of the stellar
material. In addition to this, there are the positivenegative pairs
1 C. D. Anderson, Science 76, p. 238, 1932.
* Since there is no clanger of misapprehension, we drop the subscript % of K in
the remainder of this chapter.
338
TEXTBOOK OF THERMODYNAMICS
XIX 129
created by the reaction (19.26). The strong electric attraction will
assure that the stellar matter remain electrically neutral. If the
density of positrons is z p , the total density of negatives will be
z e = z p + ZQ, the excess negative charge being neutralized by the
atomic ions. Equation (19.27) gives, therefore,
or
Z P (Z P + ZQ) = K,
z p = ((K + so 2 /4) H  20/2].
(19.30)
To fix our ideas, let us assume ZQ 5 X 10 21 cm" 3 , which is, pre
sumably, of the right order of magnitude for stars of the G and K
types. Table 49 gives the density of positrons z p for a large range of
temperatures which may or may not exist in the stars. The last col
umn lists the heat function stored up in the pairs per 1 cm 3 . From
about 5 X 10 8 degrees on, z p and x become independent of ZQ and
would be the same for all values of this density smaller than 5 X 20 21 .
At the same time the kinetic energy of the electrons becomes more
important than the intrinsic.
TABLE 49
EQUILIBRIUM OF NEGATIVE AND POSITIVE ELECTRONS
T
K
Zp
'()
\cm'/
8 X 10 7
6.0 X 10 10
1.2 X 10"
4.5 X 10~ 45
9
1.2 X 10~ 2
2.4 X 10 24
9.3 X 10 38
1 X 108
8.5 X 10 3
1.7 X 10~ 18
6.1 X 10 32
2
9.1 X 10 29
1.8 X 10 8
7.3 X 10'
3
5.0 X 10 38
1.0 X 10 17
4.2 X 10 3
5
6.2 X 10"
7.9 X 10 22
3.5 X 10 9
1 X 10"
1 . 7 X 10"
4.1 X 10 26
2.2 X 10 13
1 X 10 10
7.4 X 10'
2.7 X 10 30
5.5 X 10 17
At first sight, one is inclined to doubt if the assumption made in
section 127 is fulfilled, namely, if the index of refraction of the radia
tion is approximately 1 in matter as dense as that. However, a check
shows that such is indeed the case because the wave length which must
be here considered is that emitted in the reaction, and it is extremely
short (1.2 X 10 " 10 cm) even at low temperatures.
129. Neutrons and protons. The neutron has a mass slightly
larger than the proton. There is, therefore, the possibility of its being
XIX 129 EQUILIBRIUM INVOLVING RADIATION 339
converted into a proton under ejection of a negative electron and per
haps of a quantum of radiation. In fact, there is much reason to
believe than an intranuclear process of this type plays a role in the
emission of radioactive 0rays. Equation (19.21) takes here the form
G n G H + G e = 0,
whence v n = 1, J> H + = v e = 1, v = 1 if we characterize the neutron
and proton by the subscripts n and H +. The equation of equilibrium
(19.23) is then
The atomic weights of the neutron and proton are 1 n n = 1.0090,
1.0076, A/z = 0.0009, the (nuclear) statistical weights gn = gH+ = 2,
whence
3 4 2 v 10 9
logio K =  15.688   log 10 T  y  (19.31)
The process of dissociation is here exothermic, and, therefore, the
conditions are peculiarly inverted. The presence of excess electrons
does not depress the degree of dissociation but favors it. The equilib
rium constant does not rise monotonically with temperature but only
rises to a maximum (at T = 6.5 X 10 9 ) and then drops again. This
maximum value of K is equal to
K = 7 X 10 32 .
Of course, the last decimal figure in the atomic weights /in and
/IH+ is not quite certain and the heat of reaction is affected with a con
siderable probable error. But even if Aju were only half as large
(0.00045) the maximum value of K would be ttut little increased,
namely K = 2 X 10 ~ 30 . As the density of electrons z e in ordinary stars
(not white dwarfs) is only of the order of magnitude 10 21 (referred to
in the preceding section), we conclude that the ratio of free neutrons to
protons in them must be negligible.
This conclusion would not be safe with respect to white dwarfs.
The density in them is so enormously high that the electrons may form
a completely degenerate gas even at temperatures as high as 10 8 de
grees. The thermodynamic potential of these electrons is then (1 7.35),
<p e = fQt + jji e c 2 , and the equation of equilibrium, <p n ^>H+ <? = 0,
gives simply
1 Bonner and Brubaker, Phys. Rev. 50, p. 308, 1936.
340 TEXTBOOK OF THERMODYNAMICS XIX 130
since /B< is negligible compared with AM The molal volume of the
electrons in the white dwarfs may be as low as v = 3 X 10 ~ 6 cm 3 ,
whence, by the formula (17.29), 9 = 0.7 X 10~ 8 T. Therefore, the
electrons may be regarded as degenerate at temperatures of the order
10 s but not of the order 10 9 . In other words, the temperature within
the white dwarfs hardly rises higher than 10 9 degrees. We are, there
fore, justified in concluding that free neutrons do not play any impor
tant role in stars, including the white dwarfs.
130. Hydrogen and deuterium. As early as 1815 the English
chemist, William Prout, formulated the hypothesis that the atoms of
heavier elements are formed by the association of hydrogen atoms.
This view is confirmed by modern science and found its expression in
the very name "proton," which is derived from the Greek "TrpSros"
meaning " first " or " primary". A scientific theory of nuclear struc
ture became possible after the discovery of the neutron : every nucleus
is supposed to consist of neutrons and protons. 1 There is a consider
able body of experimental facts to support this theory. In the so
called experiments on permutation of matter it is observed how fast
moving particles knock protons and neutrons from or into atoms, con
verting them into other elements. Inasmuch as a neutron can be
formed by the association of a proton and an electron (preceding sec
tion), we can also say that any nucleus may be produced out of a
suitable number of protons and electrons, although the electrons do
not retain in it their individuality but merge with the protons, forming
neutrons.
The simplest examples of the building up of heavier elements is the
formation of deuterium and helitim, which we shall discuss in this and
the next section. There is direct experimental evidence for these
processes. When a fastmoving deuteron hits a target, phenomena
are observed which offer evidence of its being split sometimes into a
proton and a neutron. On the other hand, we have seen in the pre
ceding section that a neutron dissociates into a proton and an electron.
It must be possible, therefore, for two protons and one electron to
unite, forming a deuteron
2Gn+ + G e GD+ = 0,
PH+ = 2, v, = VD+ =1, v = 2 and
* W. Heisenberg, Zs. Phys. 77, p. 1932; 78, p. 156, 1933.
XIX 130 EQUILIBRIUM INVOLVING RADIATION 341
The atomic weight of the deuteron 1 is MD+ = 2.0142, the statistical
weight 2 is #D+ = 3. Hence A/* = 0.0015,
7 04 v 10 9
lo glo K = 35.63 + 3 logio T  ^^ (19.33)
Numerically:
TABLE 50
DEUTERON PROTON EQUILIBRIUM
T K K*
10 8 1.6 X 10 11 4 X 10~
2 X 10 8 2 X 10 26 4 X 10 12
4 X 10 8 7 X 10 48 8 X 10 21
It is clear from this equation that in the state of equilibrium there
should be very little hydrogen below 10 8 degrees. The density of H+
cannot be larger than K^ since zn+/z e is certain to be smaller than 1.
At low temperatures the conditions are somewhat different because
one cannot treat the gases as completely ionized but must regard them
as atomic or even as molecular. For atomic hydrogen and atomic
deuterium the equation would be
2G U  CD = 0,
J>H ==: 2, VD == 1 v = 1.
ZH = CKz D ) M or 2 D = Z H 2 K,
3 7 04 V 10 9
logio K = 20.260 +  lo glo T  ^~ (19.34)
Again we see that in the state of equilibrium there should be an
enormous excess of deuterium. There is little doubt that positive and
negative electrons, in our universe, are really in equilibrium in the
sense of section 127. The recombination of pairs progresses with
great rapidity, and under ordinary conditions there are no more of
them than calculated from our equations. Also, there is no direct
evidence that the free neutrons are more numerous than would be
expected theoretically. But when it comes to the formation of deu
terium from hydrogen and to the analogous processes considered in the
next two sections, we must conclude that the world is very far from
1 Bonner and Brubaker (footnote on p. 339).
2 At the high temperatures here in question there may exist and be stable another
modification of the deuteron (with zero nuclear spin). The joint statistical weight
would then be #D+ = 4 and the constant 35.50.
342 TEXTBOOK OF THERMODYNAMICS XIX 131
equilibrium, indeed. It appears that the protons, unless they are in
some unusual state of activation, are unable to combine and form
heavier nuclei.
131. Hydrogen and helium. The case of formation of helium by
the association of four hydrogen atoms is of historical interest because
it was the first example of permutation of elements to which the theory
of thermodynamical equilibrium was applied. 1 The helium nucleus or
aparticle can be formed of four protons and two electrons:
4G H + + 2G.  G a = 0,
whence *>H+ =4, v e = 2, v a = 1, v = 5, and
z H + 4 Ze 2 /z a = K or SH+ = (Kzjz^z"* .
The atomic weight of helium is HH e = 4.0040, and of the aparticle
H a = 4.0029, the statistical weight in both cases gn e = g = 1 This
gives AM = 0.0284, and
i 333 y inn
logio K = 92.44 + 7.5 lo glo T  ~  (19.35)
Numerically:
TABLE 51
HELIUMHYDROGEN EQUILIBRIUM
T K K y *
7 X 10 2 X 10 1.2X10'
8 X 10 8 4X10 8 1.4X10*
10 4 X 10 2 4.6X10'
5 X 10 3 X 10 1 ' 8 4.3 X 10"
Seeing that z a < z e and that, in ordinary stars, zj 4 is considerable
(up to 10 6 ), hydrogen should be practically absent at temperatures
below 10 9 . At low temperature, the gases must be taken as atomic
because they are no longer ionized
4G H  G H = 0,
or Z H
It is hardly necessary to calculate K in this case, because it is imme
diately obvious that the positive terms in its expression are diminished
while the negative remain the same as in (19.35) so that logio K has
an enormously large negative value. If equilibrium prevailed, with
respect to this reaction, no appreciable amount of hydrogen could
exist outside the stellar cores.
1 R. C. Tolman, J. Am. Chem. Soc. 44, p. 1902, 1922.
XIX 132 EQUILIBRIUM INVOLVING RADIATION 343
The formation of deuterium and helium are merely the beginning
of the process of association of protons into heavier elements. The
further combination of aparticles, protons, and electrons into stable
nuclei is also exothermic, taking place under a reduction of mass
(mass defect). The heavier atoms are, thermodynamically, still more
probable than helium. Although the nuclei with the largest mass
defects are the most abundant, it does not seem that their abundance
relatively to helium and to one another is in quantitative agreement
with the thermodynamical expectations. It is not certain that the
present data on atomic weights are sufficiently accurate to decide this
question. Be this as it may, it is quite certain that equilibrium does
not exist with respect to hydrogen, as has already been emphasized at
the end of the preceding section.
132. Annihilation of hydrogen. The only experimentally estab
lished case of annihilation of matter is the mutual destruction of a posi
tive and a negative electron. The annihilation of protons, hydrogen
atoms, or heavier atoms has not as yet been directly observed. How
ever, in the present state of theoretical knowledge no valid reason can
be advanced why such a particle could not completely disappear, con
verting its energy into radiation. The equations for the annihilation of
a neutron, on one hand, and of a protonelectron system, on the other,
are very little different, and we shall consider the latter because of the
greater abundance of protons. The reaction is given by
G H + + G e = 0,
"H+ = " = 1, v = 2, A/x = 1.0081, whence and from (19.25),
ZH+ = K/z*> \
i r o*c.7i r 473 X 10 12 (19.36)
logio K = 36.275 + 3 logio T
Numerically:
TABLE 52
T = 5 X 10 10 K  2 X 10"
7 X IQio 2 X 10*
8 X 10 10 7 X 10'
10 11 1 X 10"
If this process is possible, the universe is extremely far from the
state of equilibrium: no matter should exist at temperatures below
7 X 10 10 degrees.
The view was expressed that there may exist also negative protons,
unobserved, so far, because of their scarcity. If this is true, there is
344 TEXTBOOK OF THERMODYNAMICS XIX 133
the possibility of the creation and mutual annihilation of positive
negative proton pairs. The reaction is the same as in the case of pos
itivenegative electron pairs only the heat is much larger (A/* = 2.015).
It follows
logio K = 39.948 + 3 logio T 
9.48 X 10 12
(19.37)
This means that temperatures of the order of 10 12 degrees are
necessary to create such pairs. Expressed in electronvolts, the energy
of such a pair is 1.88 X 10 9 e.v.
How exceedingly small the numbers in (19.36) and (19.37) are,
appears from the following fact. A few years ago the astronomers
calculated the size of the universe as 10 87 cm 3 (in the meantime it has
become doubtful whether the universe is finite). The reciprocal of this
is still very large compared with K at T = 10 10 or at lower tempera
tures. It follows that there could not be a single proton, either isolated
or as a positivenegative pair, in the whole of the universe. The last
atom of matter should have been converted into radiation if equilib
rium prevailed.
133. Influence of gravitational fields. The interest of the proc
esses discussed in this chapter lies in their applications to problems of
cosmology. However, in treating them we have not taken into con
sideration the gravitational fields existing in and around the stars. It
will be well to say a few words of explanation why the effects of the
gravitational fields can be neglected on account of their smallness.
We have treated the problem of equilibrium of heavy gases in a gravi
tational field in section 108 and have seen there that the only correc
tion which the equations need is a term
added to the equilibrium constant log K. This constant, as given by
eq. (19.23), already contains the term A/ic 2 /jRr, which, combined
with the correction for gravitation, becomes
(1 + Q g /c 2 )&'C 2 /Rr.
In other words, the heat of reaction or the energy available in the
process is changed in the proportion (1 + & g /c 2 ) owing to gravitational
forces:
V = (1 + fl,/c 2 )A/u. (19.38)
XIX 133 EQUILIBRIUM INVOLVING RADIATION 345
The gravitational potential Q g vanishes at a great distance from a
star and is negative in its vicinity, having its minimum at the surface.
The factor is, therefore, smaller than 1, meaning that the heat of
reaction is decreased. But even the largest known gravitational
potentials, those of the white dwarfs, are materially smaller (in abso
lute value) than 0.002 X c 2 , whereas within the accuracy of the preced
ing chapter a correction of 0.2% is negligible.
Equation (19.38) is, of course, quite general and applies to any
molecular or atomic process in which energy is converted into radia
tion. According to the quantum theory, the frequency v of the
emitted radiation is connected with the available energy per atom or
molecule AE = c*&n/n A by the relation AE = hv, whence
v' = (1 + tt g /c 2 )v. (19.39)
The gravitational field changes the frequency of radiation emitted
in an elementary process, decreasing it (gravitational red shift), a
phenomenon predicted by Einstein and since confirmed experimentally.
We have based these remarks on the Newtonian theory of gravita
tion. On general grounds we are justified in saying that the condi
tions cannot be materially different in the general relativity of Ein
stein as long as tt g /c 2 is small. In fact, chemical equilibrium was
investigated from the standpoint of the general relativity by Tolman, 1
who found the same relations as in the absence of any gravitational
fields. Einstein's formula of the gravitational red shift
/ = [(1 + 212,A 2 )
is, in practice (!2 g /c 2 1), identical with (19.39).
1 R. C. Tolman, Proc. Nat. Acad. Sci. 17, p. 159, 1931.
CHAPTER XX
MAGNETIC AND ELECTRIC PHENOMENA
134. Langevin's theory of magnetization. The systems treated in
the preceding chapters were completely described by the thermo
dynamical variables p, V, T, apart from data about their composition.
Even in the case of charged gases, discussed in sections 106115, the
electric potential was regarded as only a part of the energy constant
and not as an additional parameter of the system. It is instructive
to consider also variables of a nonmechanical nature, and we are going
to do so in this chapter, beginning with the theory of magnetics or
substances capable of magnetization. It will be sufficient to treat the
case of isotropic substances because the generalization is obvious and
would only encumber our expressions without adding anything essen
tial to them. Suppose that a system of this sort is placed in a homo
geneous magnetic field whose strength is H in the absence of the system
(i.e. before the system is brought in). Let M denote the component
of the total magnetic moment (or total magnetization) in the direction
of the field.
We know from the theory of electromagnetism that the work which
the field does, in raising the magnetization from M to M + dM, is
DW' = HdM. In accordance with the convention of section 7, we
count the work done by outer forces against the system as negative.
The total element of work (2.02), including the mechanical and mag
netic parts, is then
pdV  HdM
and eq. (3.04) of the first law has the form
DQ  dU + DW = dU + pdV  HdM. (20.01)
The state of the magnetized system depends on the two thermal
parameters T, V we used all along and, in addition, on the magnetic
parameter M. In order to have a complete description of the system,
we must know the three functions
U = C7(r, F, M), p p(T, F, M), (20.02)
H  H(T, V, M) (20.03)
346
XX 134 MAGNETIC AND ELECTRIC PHENOMENA 347
In the sense of section 10, we may call these three relations the
caloric, the thermal, and the magnetic equations of state. Of course,
it is arbitrary which of the six parameters we regard as independent.
For most of the applications it will be convenient to describe the sys
tem in terms of T, p, H, but occasionally we shall use also T, p, M.
The first application of the second law of thermodynamics to such
a system was due to Langevin. 1 As he pointed out, in the expression
(4.07) of the entropy differential, dS = DQ/T, the magnetic work
gives rise to the additional term
DW II
dS' = = dM, (20.04)
and he postulated that dS' be an exact differential. This is possible
only when the factor H/T is a function of M (compare section 8) and,
vice versa, M is a function of H/T,
/TT\
(20.05)
This is the form of the magnetic equation of state (20.03), accord
ing to Langevin. It is quite obvious that this law is not general, a
case in point being the behavior of diamagnetic substances whose mag
netization is known to be independent of temperature:
M =  AH, (20.06)
where A is a constant. 2 What is, then, the special assumption made
in this theory and leading to the formula (20.05)? It lies in the sep
arate treatment of the magnetic part of the entropy (20.04) as an
exact differential, independently of the remaining part
(dU + pdV)/T. (20.07)
This implies, of course, that the expression (20.07) is also an exact
differential which cannot depend on the variable M (or H) since its
differential is absent in it. Consequently the functions U and p must
be also independent of M. Instead of (20.02),
U = U(T, V), p = p(T, V). (20.08)
We met with a similar case in the theory of perfect gases. In the
expression (4.16) of dS the two terms are separately exact differentials,
1 P. Langevin, Ann. Chim. Physique 5, p. 70, 1905.
8 We speak here of the ordinary or atomic diamagnetism. There exist a few
substances (as bismuth, antimony, etc.) with socalled crystal diamagnetism which
has a temperature coefficient.
348 TEXTBOOK OF THERMODYNAMICS XX 134
and this is due to the internal energy depending only on the thermal
variable T. By analogy we shall call a substance whose equations of
state follow the simple laws (20.05), (20.08) a perfect magnetic.
Langevin initiated also the statistical theory of magnetization and
gave an explicit expression of the law (20.05) for paramagnetic mate
rials. As corrected to take into account the restrictions of the quan
tum theory, it can be written
M 
(20.09)
and is sometimes called the LangevinBrillouin formula. Z represents
here the total number of atoms, and the product jgfo the total magnetic
moment of one atom. The factors have the following meaning:
o the Bohr magneton
A) = ^ = (0.9174 db 0.0013) X 10 20 erg gauss 1 , (20.10)
g the Lande factor
3
* 2+ 2/C7+1) ' ( }
j is the quantum number of the total angular momentum of the atom,
compounded of the resultant spin momentum s and the resultant
angular momentum /. At 20 C the numerical value of fo/kT is
2.28 X 10~ 7 : since jg is rarely larger than 7, the argument of the
function F is small (< 0.050) at room temperature, even for fields of
the order of 30 000 gauss. Under these circumstances all terms of the
expansion of F(a) into a power series are negligible, excepting the first
linear term, and the magnetization is well represented by Curie's law
*, CH
M = , (20.12)
where C is a constant. For large values of the argument, the function
F asymptotically approaches the value 1 (magnetic saturation), but its
indications appear only at very low temperatures.
The formula (20.09) was derived for gases and solutions of salts
in which the atoms or ions can freely rotate. Even for these sub
stances its validity is not general, being often impaired by doublet or
multiplet structure of the energy levels. Only when the width of
XX 135 MAGNETIC AND ELECTRIC PHENOMENA 349
these multiplets is very small or very large compared with kT is the
formula applicable. It is remarkable, however, that it represents
with great accuracy the magnetization of solid paramagnetic salts of
rare earths, especially those in which the paramagnetic atoms are not
very close together but separated by nonmagnetic atoms (as, for
instance, in hydrates). The classical example, in this respect, is
hydrated gadolinium sulfate investigated in Leiden. 1 The magnetic
moment of this substance was measured down to 1?3 K and found to
obey strictly the law (20.09) with j = J and g = 2. At the lowest
temperature and highest field the saturation was far advanced, the
magnetization being 95% of the possible maximum.
Exercise 106. Calculate the entropy term S' from (20.04) for substances whose
magnetic moments are expressed by (20.09). Show that
S' = kZ\ log sinh ( a ~\ ) /sinh  aF(a) \ f const.
I L \ 2j/f 2/J J
Exercise 107. In his original formula Langcvin assumed that the orientations
of the molecular magnetic moments are not restricted by quantum conditions.
This formula can be obtained from (20.09) by the following transition to the limit:
let j go to oo, while goes to zero in such a way that jg&o = /3 remains constant
where j3 represents the molecular magnetic moment). Show that the result is
M = Z/5 f(0H/kT), f(a) coth a  I/a.
135. Magnetothermal and magnetocaloric relations. It will be
well to derive a few relations which will elucidate still further the
thermodynamical status of perfect magnetics and diamagnetics. We
start from the generalized thermodynamic potential
* = U  TS + pV  HM, (20.13)
whose differential is, according to (20.01) and (4.07),
d& =  SdT + Vdp  MAIL (20.14)
Hence
, V = ( J , M =  P^j (20.15)
Substituting this into (20.13)
*  < 20  16)
\dP/T,H \d
1 H. R. Woltjer and H. Kamerlingh Onnes, Leiden Comm. 167 b, c.
350 TEXTBOOK OF THERMODYNAMICS XX 135
The three partials (20.15) give reciprocity relations when differ*
entiated a second time. Thus we obtain from the second and third
p.T T.H
This is a relation between measurable quantities. The left side
is the change of volume due to the field, its relative value (divided
by V) is called volume magnetostriction or Barrett effect. The right
side represents the change of the total magnetic moment due to pres
sure. Such an effect was foreseen and first measured by Nagaoka and
Honda 1 and is named after them. Although accurate modern mea
surements of both effects are available, they were not carried out under
comparable conditions and are not suitable for a quantitative test of
the relation (20.17). It should be noted that in perfect magnetics
the volume does not depend on the magnetic parameters (preceding
section). The experimental existence of magnetostriction in ferro
magnetic materials shows, therefore, that these substances do not
strictly follow Langevin's law (20.05). However, the effect is very
small and the deviation, presumably, not large.
Exercise 108. In a particular specimen of iron (at / = 20 C, H = 1 gauss)
the partial (dM/dp)T,H was found to be 3 X 10 ~~ 3 gauss per unit volume and per
1 atm. Calculate the volume magnetostriction from formula (20.17). (Use
absolute units).
The other two reciprocity relations following from (20.15) are
no
97 / ptH
We shall use them for transforming eq. (20.16) as follows. We
differentiate (20.13) partially with respect to p, and with respect to H,
and take into account the relations (20.17) and (20.18):
(
\
In the case of perfect magnetics we have (dU/QH) TtP
0, and from (20.17) and (20.20)
1 Nagaoka and Honda, Phil. Mag. 46, p. 261, 1898; see also S. R. Williams,
Int. Crit. Tables VI, p. 439, 1929.
XX 135 MAGNETIC AND ELECTRIC PHENOMENA 351
This is only another proof of Langevin's result because the general
integral of this equation is the expression (20.05). On the other hand,
the law (20.06) of diamagnetism gives OJ7/9#)r, P = ~~ AH = W> so
that the internal energy contains a magnetic term whose differential is
dU' = (dU/QH) Ttp dH = MdH = HdM. (20.21)
A look at the structure of the element of heat (20.01) shows that
the additional term in dU just cancels the work of the magnetic field
DW' = HdM, so that DQ is independent of magnetic parameters.
We see from this that diamagnetic and perfect magnetic substances are
complete opposites with respect to the caloric action of putting them
into a magnetic field. In diamagnetics, the (positive) work against
the field is done at the expense of the internal energy without any
contribution from external heat sources. It is a process at the same
time adiabatic and isothermal: the loss is sustained, as it were, by the
constant of the internal energy and does not affect the temperature of
the body or its equilibrium with the environment. On the contrary,
the (negative) work of magnetizing perfect magnetics is completely
supplied by outer sources, when conducted isothermally. When the
process is adiabatic it leads to a rise of temperature in the sub
stance.
In the interest of the following section we shall say here a few
words about the heat capacity of magnetics. In treating simple sys
tems, we distinguished in section 14 the specific heats at constant vol
ume c v and at constant pressure c p . Similarly, we have to speak here
about the following two heat capacities. The heat capacity at constant
magnetization (and, say, constant pressure) is the heat that must be
imparted to the magnetic when its temperature is raised by 1 degree
and, at the same time, the magnetic field is changed in such a way as
to keep its magnetization constant: C P M = lim(AQ/Ar) p Af. According
to the expression (20.01), no part of this heat is used for magnetic
work. On the other hand, the heat capacity at constant field strength
C P H = Hm (&Q/&T) P H refers to a process accompanied by a change of
M and involving magnetic work. In view of the relation AQ = T&S,
we can also write
The relation between these quantities analogous to (4.27) is found,
in the simplest way, by considering 5 a function of T and M, while M ,
in its turn, depends on T and H, i.e. S = S[T, M(T, H)]. (We need
352 TEXTBOOK OF THERMODYNAMICS XX 136
not bother about the pressure p, since it is assumed to be always con
stant.) The rules of partial differentiation give us
,
dT/ P> M \dM/ PtT \d ptH
In the partial (dS/dM) PtT the variables p and T are considered as
constant so that M is a function of H only: therefore,
Substituting this and making use of eq. (20.18), we find the desired
relation
r c r
C vn  C PM = T 
136. Cooling by adiabatic demagnetization. Adiabatic demag
netization of a paramagnetic substance is analogous to the adiabatic
expansion of a simple system. In either case, the work against external
forces is done at the expense of the internal energy of the system and
leads, in general, to a decrease of its temperature. The use of this
" magnetocaloric effect " for the production of very low temperatures
was proposed independently by Giauque J and Debye. 2 The usual
method of lowering the boiling pressure of liquid helium does not per
mit in practice 3 to reach temperatures below 0.7 K. Work with
the magnetocaloric method has been under way in recent years, at
Berkeley (California) under the direction of Giauque, 4 and at Leiden
(Holland), under the direction of deHaas. 5 It has been very suc
cessful in bringing the range of accessible temperatures considerably
nearer to the absolute zero point.
The equation of the adiabatic dS = dQ/T = is, according to
(20.01),
dS = (dU + pdV  HdM)/T = 0, (20.24)
when S is expressed as a function of T, p t H, also
(
a\
<3 T/ VtH \ OP/ H t T ^O** / T.p
i W. F. Giauque, J. Am. Chem. Soc. 49, pp. 1864, 1870, 1927.
* P. Debye, Ann. Physik 81, p. 1154, 1926.
Keesom, Proc. Amsterdam 35, 136, 1932.
W. F. Giauque and C. W. Clarke, J. Am. Chem. Soc. 54, p. 3135, 1932;
W. F. Giauque and D. P. McDougall, Phys. Rev. 43, p. 768; 44, p. 235, 1933.
W. J. deHaas, E. C. Wiersma, and H. A. Kramers, Physica 13, p. 171; 1, p. 1,
1933; W. J. deHaas and E. C. Wiersma, Physica 1, p. 1107, 1933.
XX 136 MAGNETIC AND ELECTRIC PHENOMENA 353
Substituting from (20.22) and (20.18),
c f(fr) *+(!!?) w  < 20  25 >
i \0l/p,H \0* / Pt ff
The procedure consists in precooling the specimen in vacuo, in a
strong magnetic field, and then turning off the field.
With p = 0, eq. (20.25) gives for the rate of the adiabatic cooling
or from (12.23)
dH
As was pointed out in the preceding section, the heat capacity C P M
does not contain any magnetic work. Therefore, it should not be
materially different from the ordinary capacity C p as measured in the
absence of a magnetic field. Of particular interest are the perfect
magnetics: from the kinetic point of view, they are the substances in
which every atom is free to adjust itself in the field without being
hindered by its neighbors. Other things being equal, they should
show the highest magnetization and the highest caloric action. For
this reason, both Giauque and Debye suggested the use of gadolinium
sulfate as a testing substance because it was known to follow the law
(20.09) down to the lowest temperatures. If we denote x = H/T,
this law implies
dM\ __x_dM /aM\ 1 dM
dT/ Pt a~ ~~ T doc ' \9/// Pl r"" T dx '
\dT/ Pi H J. ax \ou/p,r
so that (20.27) becomes
dT _ x(dM/dx)
dH~ C pM +x 2 (dM/dx)'
(20.28)
Of course, in the case of perfect magnetics there is no need to use a
differential formula because the two terms (with p = 0), in the formula
(20.24), are separately integrable. At zero pressure there is no differ
ence between c p and c v , and the energy differential is dU CvdT.
Hence
(20.29)
354 TEXTBOOK OF THERMODYNAMICS XX 137
If the specific heat follows the third power law (18.70), the first
integral is equal to 6.45 X 10 9 (r/9) 3 erg/deg mol. Suppose we start
the demagnetization at the temperature To and the field HQ. When
the field is reduced to H = 0, the temperature has the value
s*XQ J
To 3  1.55X10 10 8 3 / x
(20.30)
where XQ = Ho/To and m is the molal magnetic moment.
It was found, however, that the specific heat of gadolinium sulfate
does not follow the third power law but exhibits an anomaly due to
multiplet structure of the lowest energy level (compare section 117).
The curve reproduced on p. 327 was obtained by Giauque and McDou
gall by measuring the magnetocaloric effect dT/dH and by calculating
C pM from a formula equivalent to (20.27). Although they reached a
temperature of 0?287 K, this substance is not particularly suitable for
magnetocaloric cooling. At the time of the writing, the greatest
success was obtained by deHaas with a mixture of two alums,
K 2 SO 4 Cr 2 (SO4)3 24H 2 O + 14.4K 2 SO4A1 2 (SO4)3 24H 2 O. Demagnet
izing from 24 075 gauss to 1 gauss, he reached 0?0044 K. In all this
work the measurements themselves were used to establish the scale
of absolute temperatures by the method explained in section 29.
However, the last figure (0?0044 K) rests, in part, on an extrapolation
and has only approximate validity.
Exercise 109. Consider a substance satisfying eq. (20.30) whose magnetization
obeys the law (20.09), with j = , g = 2. Start with T 3 K and H = 25 000
gauss and demagnetize to H = 0. What will be the final temperature if = 200?
137. Supraconductivity and thermodynamics. It was discovered
by Kamerlingh Onnes that the electric resistance of certain metals
suddenly drops to practically nothing when they are cooled below tem
peratures called their transition points. We give the transition points
of the principal supraconductive substances in Table 53.
The transition point is displaced when the system is put into a
magnetic field H. As the magnetic field is increased, the transition
point moves continuously to lower temperatures, and at a certain
strength of it (threshold value) the supraconductive state ceases to
exist altogether. From the point of view of electromagnetism, the
immediate effect of switching on a field consists in setting up currents
in the supraconductor, in cpnformity with Faraday's law of induction.
But the conductivity of these substances is so large that the currents
continue to flow, without apparent loss, as long as the state of supra
XX 137
MAGNETIC AND ELECTRIC PHENOMENA
35S
TABLE 53
SUPRACONDUCTIVE TRANSITION POINTS
Elements
Elements
Intermetallic
compounds
Intermetallic
compounds
MB
70 K
V 4 3 K
Au a Bi 1 84 K
ZrB 2 82 K
7n
78
Nb 9.2
CuS... 1.6
TaSi . . . . 42
CH
0.6 (ca)
Ta 4.4
VN.... 1.3
PbS .... 41
HfT
4 22
La 47
WC 2 8
**is
Al
1 14.
WaC 2 05
Ga
. . 1 05
Intermetallic
MoC. 7 7
Alloys
Tl
9 37
A/fnP 9 1
Ti
1 75
TIN 1 4
PbSnBi 8 5
Th
1 5
BiflTlj. 6 5
TiC. 1 1
PbAs. 8 4
Sn
3.71
Sb 2 Tl 7 .... 5.5
TaC... 9.2
PbAsBi 9.0
Ph
7.2
Na 2 Pb 6 ... 7.2
NbC... 10.1
PbBiSb 8.9
In
3.37
Hg 6 Tl 7 ... 3.8
conductivity is maintained: they are therefore called persistent cur
rents. Another result also follows from the theory of electromagnet
ism: when the magnetic field is switched on, after the supraconductive
state is established, neither the field nor the currents can penetrate to
any appreciable depth of the metal. The persistent currents flow at
the surface and produce a magnetic field which exactly compensates
within the metal the external field. Experimental work has shown
that the same conditions prevail when the metal is cooled below the
transition point in an already existing magnetic field. The lines of
magnetic induction are pushed out of the metal, as it becomes supra
conductive, until the induction in the interior vanishes. This result
was first announced by Meissner and Ochsenfeld 1 and has been con
firmed since by many investigators. Although there is still some
doubt whether it holds rigorously, it describes the phenomena with an
accuracy sufficient for thermodynamical purposes. In short, a supra
conductor behaves in a magnetic field like a substance of the perme
ability /i = 0. Since /* is connected with the susceptibility K by the
relation /* = 1 + 4?r/c, the supraconductors can be formally described
as diamagnetics with the susceptibility K = 1/47T.
Let us consider a long stretched (needleshaped) supraconductive
body placed with its axis in the direction of the magnetic field. Accord
ing to the electromagnetic theory, the outer magnetic field produced
by a body of this shape is negligible . Its j(negative) magnetic energy
1 W. Meiasner and R. Oc'senfeld, Die Naturwiseenachaften 21, p. 787, 1933.
3S6 TEXTBOOK OF THERMODYNAMICS XX 137
is due to the absence of any field in the interior and is equal to
per unit volume, or to
(20.31)
per mol. The magnetization (per unit volume) is an expression of the
type (20.06) characteristic of diamagnetics. We can apply, therefore,
to supraconductors the results established in the preceding section for
diamagnetic substances: the internal energy contains U H as an additive
term, while the entropy is independent of the magnetic parameters.
Consequently the thermodynamic potential (<p = u Ts + pv) has,
in the case of supraconductive materials, the form
vH 2
**> , (20.32)
O7T
where w is the potential in the absence of a field. At the very low
temperatures we are here considering, the pressure effects are
negligible: there is no observable change of specific volume, so that
it is permissible to ignore the pressure as a variable and to regard
<po(T) as a function of temperature only, and v as a constant. The
normal (nonsupraconductive) modification of the metal is, as a rule,
nonmagnetic, so that its thermodynamic potential does not contain
any magnetic term but is, simply, <p n (T). The equation of equilibrium
between the two phases becomes, therefore, A3> = ^> n <p = or
A* = Vn (T)  ,(D + ~ = 0. (20.33)
O7T
Of course, it is not a priori certain that the supraconductive state
is in true equilibrium and that the second law can be applied to it.
The hypothesis that it represents a phase in the thermodynamical
sense was first made by Langevin. 1 It was corroborated by the great
sharpness of the transition point in good crystalline specimens. The
experimental test of the equation (20.33) is also reassuring: its conse
quences hold with good accuracy. This relation gives the dependence
of the transition point on the strength of field H and can be tested in
several ways. The first and second partials of A$ with respect to T
are, according to (5.37) and (7.22),
1 P. Langevin, Rapports du 1* Conseil & vay, p. 301, 1911.
XX 137
MAGNETIC AND ELECTRIC PHENOMENA
357
The differentiation of (20.33) gives the equation (l/T)dT
(vH/v)dH = or
analogous to the ClausiusClapeyron equation. It was tested by
Keesom and Kok l and found to be in agreement with experiments.
Of great interest is the observation of these authors that the latent
heat decreased as the field diminished and could not be measured at
all in the case of field free transitions (i.e. dH/dT 7* oo for H = 0). If
this observation is confirmed, it would mean that the equilibrium, in
the case H = 0, is of the second order (section 49). Let us consider
this case: we denote the transition point in the absence of a field by
To and express A$, for T TQ + dT, H = dll, to terms of the second
order. From (20.34)
l/Ao\
2\r /
(dry  
or
AO _ v (dlf^
TQ 4ir\a
(20.36)
This equation was obtained by Rutgers 2 and in a different way by
Gorter and Casimir. 3 All these authors regarded the equilibrium in
question as one of the second order. The equation (20.36) was con
firmed with remarkable accuracy by the beautiful measurements of
Keesom and Kok on tin and thallium. 4
TABLE 54
ToK
v
cm l mol" 1
dH/dT
gauss deg~ l
Aco (calc)
cal deg^mol"" 1
Aco (obs)
Tin
3.71
14 2
151 2
00229
0024
Thallium
2.36
16 9
137 4
00144
00148
Nevertheless we do not agree with Rutgers, Gorton, and Casimir.
1 W. H. Keesom and J. A. Kok, Physica 1, pp. 503, 595, 1934.
1 Appendix to Ehrenfest's paper, note on p. 128.
* C. J. Gorter and H. Casimir, Physica 1, p. 306, 1934.
4 W. H. Keesom and J. A. Kok, Physica 1, p. 175, 1933.
358
TEXTBOOK OF THERMODYNAMICS
XX 137
The equilibrium cannot be of the second order since an unsurmount
able difficulty is introduced by the double sign of
dH ^ ATT A^oV
dT " \v To/
Represented graphically (Fig. 60), the transformation is of the
general type of Fig. 18 on p. 132. With any plausible choice of the
functions <f>o(T), <p n (T), the analysis shows that the lines representing
the equilibrium continue beyond the point TV Therefore, both
regions labeled (1) in Fig. 60 correspond to the supraconductive state
so that cooling along the line // = does not involve any transition.
We maintain, therefore, that the equilibrium in question is, in reality,
FIG. 60 FIG. 61
Supraconductive transition point in its dependence on magnetic field.
of the first order, there being a small but finite latent heat. According
to the general discussion of section 49, this would imply a transition
curve of the shape given in Fig. 61. That Rutgers' formula is satisfied
with such a remarkable accuracy is due to the following coincidence.
Equation (20.35) holds for the whole curve and can be differentiated
along it. Since // T = As, its differential is (Ac/T)dT, and we find
Ac
T
(20.37)
Where the line of Fig. 61 is straight (d 2 H/dT 2 = 0), the slope is given
exactly by Rutgers* expression, whether the equilibrium in TO is of the
first or of the second order. 1
There were attempts to make other aspects of supraconductivity
amenable to a thermodynamical treatment. However, they failed to
1 In a later paper (Phys. Zs. 35, p. 963, 1934) Gorter and Casimir admit that the
supraconductive state would be thermodynamically stable above the transition
point "if it existed." They postulate, therefore, that it cannot exist. This seems
an adhoc hypothesis, and the difficulty is much better disposed of by the simple
assumption made in the text, that the equilibrium is of the first order.
XX 138 MAGNETIC AND ELECTRIC PHENOMENA 359
give a complete picture of the phenomena for which they were intended.
Limitations of space do not permit us to enter into their discussion.
138. Electrostatic phenomena. Dielectrics or substances capable
of electric polarization P form, in many respects, an analogue to the
magnetics treated in section 134. The work which must be done on
a dielectric body in order to increase its total electric moment (or
polarization), in the direction of the field , from P to P+ d P, is
E dP,
so that the complete element of work becomes
DW = pdV  E dP, (20.38)
in analogy with the magnetic expression of section 134. There are
two types of dielectric polarization which, from a thermodynamical
point of view, correspond closely to the types of magnetization in
atomic diamagnetics and in perfect magnetics.
(A) The Lorentz type of polarization. If the molecules of a sub
stance have no electric moments in their normal state, the application
of an electric field creates such moments in them by displacing the
electrons from their normal position. The expression for this sort of
polarization (taking account of the field of the electric moments) is
P AE', (20.39)
where A is a constant and E = E + P is the effective field strength.
o
The phenomenon is mainly intramolecular, and A is independent of
temperature. This has the thermodynamical implications discussed
in connection with diamagnetism.
(B) Dipole polarization. It was pointed out by Debye l that some
dielectric substances may possess permanent molecular dipoles. He
suggested, therefore, taking over Langevin's formula (developed for
magnetic dipoles) in order to account for empirically observed changes
of electric polarization with temperature. The case of polarization is
simpler in two respects. In the first place, the electric dipoles have
no tendency to orient themselves in discrete quantized directions.
The original Langevin formula derived on lines of classical statistics
holds for them: this means that one has to take in eq. (20.09) j = oo
and jgPo = p e = const. 2 In the second place, the conditions in dielec
1 P. Debye, Phys. Zs. 13, p. 97, 1912.
2 The transition to the limit is the same as in exercise 107. When the argument
a is small, there holds the expansion coth a I/a f Ja; therefore, the function of
exercise 107 becomes /(a) = ^a, leading to formula (20.40).
360 TEXTBOOK OF THERMODYNAMICS XX 138
tries hardly ever approach saturation so that the term of first order in
E! IT represents an excellent approximation
.
where fa is the electric dipole of a molecule and Z the total number of
molecules. Of course, one should expect this expression to hold only
in substances where the dipoles can freely rotate, i.e. in gases and
liquids. In analogy with the perfect magnetics of section 134 they may
be called perfect dielectrics.
In general the dipole dielectrics possess also an appreciable polariza
tion of the Lorentz type so that very few of them are, even approxi
mately, perfect. The dielectric constant D (which is accessible to
direct measurement) is connected with the polarization by the rela
tion (D 1)/(D + 2) = $irP/VE' and has the general expression
F+I" iv + i z kT' (20  41)
with z = Z/F. By plotting D against \/T one can separate out the
part due to dipoles and determine the moment. This proved a very
helpful method for investigating the properties of many organic
molecules. 1
The electric analogue to magnetostriction is the phenomenon of
electrostriction. Because of the close parallelism of the thermody
namical properties of dielectrics and magnetics we can directly take
over eq. (20.17), writing
T.B
We see that electrostriction (represented by the left side) is asso
ciated with another effect called piezoelectricity, which consists in the
dependence of the polarization P on pressure. Our formula refers to
the volume effect, but piezoelectricity is mostly observed in crystals,
where it is particularly strong in certain crystallographic directions.
In turmaline the piezoelectric effect (increase of a component of the
polarization per unit volume and per unit increase of a strain com
ponent) is 5.78 X 10~ 8 in e.s. cgs units, in quartz 6.9 X lO" 8 , but
it is particularly large in Rochelle salt, where it reaches 8100 X 10"" 8 .
An electromagnetic wave, acting upon a plate cut from a piezoelectric
crystal, periodically changes its size, because of the electrostriction
' See P. Debye, Polare Molekeln, Leipzig, 1929.
XX 139 MAGNETIC AND ELECTRIC PHENOMENA 361
associated with piezoelectricity, and sets up in it elastic oscillations.
The amplitude is particularly large when the frequency of the wave is
in resonance with the characteristic period of the elastic oscillations
in the plate. Such piezoelectric resonators have received important
technical applications in the last decades. According to eqs. (20.02),
the deeper reason for the existence of piezoelectricity is the dependence
of the pressure (or strain) on the polarization P. It is to be expected
that the internal energy and entropy should then also depend on P
and H. From (20.18) we can, therefore, conclude that piezoelectric
substances are at the same time pyroelectric, i.e. their electric moment
is influenced by a change of temperature. 1
It is here the place to make a few supplementary remarks about
the theory of strong electrolytes (section 115) and to explain why we
could not calculate the internal energy U directly but had to take
the indirect way over the work function \I>. The reason is that the
electric or magnetic energy of a system cannot always be regarded as
a part of its internal energy. We have seen that such simple condi
tions obtain only in the case of atomic diamagnetism or of polarization
of the Lorentz type, when the electromagnetic work does not depend on
the thermal parameters. The opposite extreme are the perfect mag
netics and dielectrics in which U remains completely unaffected by
the electromagnetic energy of the system. The strong electrolytes
occupy an intermediate position. The addition U to the internal
energy due to the ionic interaction should be calculated from the
expression (17.64) of the work function V e by means of the relation
. = U e  TS e or
TJ ty _ T\  
u * ~~ ^* l \ IT
\9* / v
The difficulty is, however, that * e depends on T not only explicitly
but also implicitly through the dielectric constant D. In a similar
way, the thermodynamic potential $ is connected with *, by the
formula $ ^$% + P*V, or
and here again the dependence of D on the volume is not sufficiently
well known for a quantitative evaluation.
139. Thermoelectric phenomena. In section 109 we introduced
the concept of the potential difference built up by thermoelectric action
1 For a discussion of these phenomena see: Geiger and Scheel, Handbuch der
Physik, Vol. XIII. Berlin 1928.
362 TEXTBOOK OF THERMODYNAMICS XX 139
across a junction of two conductors. The discovery of the e.m.f.
(electromotive force) of thermoelectricity was made by Seebeck in
1822. It was supplemented in 1834 by observations of Peltier's on a
peculiar heat development due to electric currents flowing through
thermoelectric junctions. Let / denote the current flowing from con
ductor (1) to conductor (2): in order to maintain the junction at con
stant temperature, it is necessary to impart to it, in unit time, the
heat
(20.42)
where Hi2 is the socalled Peltier coefficient, depending on the nature of
the conductors and on the temperature of the junction. IIi2 is positive
(i.e. heat must be imparted to the system) when the current / has the
same direction as that produced by the thermoelectric action of the
junction. The sign of the Peltier coefficient is reversed when the cur
rent flows in the opposite direction (1X21= IIi2). This fact gives
the experimental possibility of separating it from Joule's heat, which
does not depend on the direction of the current. From the theoretical
point of view, the same fact is taken as an indication that the develop
ment of Peltier heat is a reversible process, amenable to thermodynami
cal treatment.
The final step was the discovery (1854) by William Thomson, later
Lord Kelvin, of the following phenomenon: when the current / flows
through a wire of homogeneous material and crosssection, but of non
uniform temperature, heat must be supplied in order to maintain the
temperature gradient. To an element of the wire with the tempera
ture rise dT, there must be imparted in unit time the heat
dQ = <rdTJ. (20.43)
a is called the Thomson coefficient, and it is positive when the current
flows in the direction of rising temperature. The Thomson effect is
reversible, in the sense that dQ changes its sign when the direction of
the current is reversed.
The thermodynamical theory of thermoelectricity is due to Lord
Kelvin. Let us consider the circuit of Fig. 62, consisting of two
wires (1) and (2) whose junctions are kept at the respective tempera
tures T and T'. When the current J is sent through the (closed) cir
cuit, the total heat which must be imparted to it in unit time to keep
the conditions stationary is, according to (20.42) and (20.43),
(20.44)
XX 139 MAGNETIC AND ELECTRIC PHENOMENA 363
From the point of view of the first law of thermodynamics, Q is the
energy which maintains the circulation of the current. In the theory
of electricity this energy has the expression
Q  /, (20.45)
where E is the e.m.f. (electromotive force) of the circuit, J being the
electric charge flowing through any crosssection in unit time. Equa
tion (20.44) indicates how this energy is spent. There is a profound
analogy between the conduction electrons streaming across a thermo
j unction and a gas forced through a cotton plug in the JouleThomson
process (section 15): the energy which the electron gas gains or loses
is equal to the difference of its heat functions in the two conductors,
Q = AX. The same is true for the Thomson effect (20.43): dQ = dX.
If the molal heat function is denoted by x and the molal electron charge
by/ = F = n A e, eqs. (20.42) and (20.43) can be written
_ (X2  Xl) Idx /IA,HC\
H 12 =  j , * (20.46)
Combining (20.44) and (20.45), we can write
/T'
((72  ai)dT, (20.47)
*
as the expression of the first law of thermodynamics. In particular,
when the temperature difference of the junctions is infinitesimal,
T  r = dT
gf? + * (20.48,
Further relations can be obtained from the second law, assuming
that the heat items of eq. (20.44) are imparted to the system in a
reversible way and that the thermoelectric effect satisfies the entropy
principle separately. Let us follow a portion of the electron gas,
responsible for the current, in its motion around the circuit back to
the initial position. The process is cyclic in that the final state is
identical with the initial and the total entropy change in it must be
zero: 2 AS = SAQ/r. The items of heat A() imparted to the electron
gas are the same as in eq. (20.44). We find, therefore,
(20.49)
364 TEXTBOOK OF THERMODYNAMICS XX 139
or taking again T T'  dT,
02
dT
The ratio Uvt/T is called the thermoelectric power of the pair of
conductors. There follows from (20.48)
dE
""" rp 9
(20.50)
* 2 ffl T dT\Tj * dT 2 '
Finally, the third law of thermodynamics has also a bearing on
thermoelectricity. Since Iliz/T represents the change of entropy
which the electrons (or other carriers) undergo in passing across the
junction, Nernst's postulate (13.07) gives directly
lim r . (H/r)  0. (20.51)
Hence the integrated form of (20.50) becomes
/ T (7 ~ (T
2 ^ * dT. (20.52)
1
This equation implies, of course, that the integral must be con
vergent, so that lim r . (^2 0*1) = 0. But not only the difference
must vanish (in the limit T = 0), but also <n and a* individually. In
fact, we see from (20.46) that <r has the physical meaning of the specific
heat of the carriers: eq. (13.15) requires, therefore,
lim r ..o <r *= 0. (20.53)
The empirical representation of the thermoelectric e.m.f. is usually
given by the equation
E  (at + 10 2 JW + 10* $d] X 10 volt, (20.54)
supposing that one of the junctions is at C, the other at f C. A
few data are given in Table 55.
From (20.54) we find (neglecting c)
a  (II/TVo, **  n  bT.
It is interesting to note that the coefficients a and b in some cases
have opposite signs. Because of the parabolic character of the
formula (20.54), as the temperature is increased the e.m.f. rises to a
XX 139
MAGNETIC AND ELECTRIC PHENOMENA
365
maximum, then declines again, and finally reverses its sign. For
instance, in the case of CdPb the reversal takes place at about
300 C.
TABLE 55
THERMOELECTRIC E.M.F.
Range in C
(1)
(2)
a
b
c
From
To
Ag
Pb
200
3.34
0.85
Al
Pb
200
0.50
0.17
Au
Pb
200
2.90
0.93
Bi
Pb
100
74.4
3.2
Cd
Pb
200
2.62
1.79
Cu
Pb
100
2.76
1.22
Fe
Pb
230
100
16.65
2.97
26.75
Ni
Pb
200
19.07
3.02
Pt
Pb
100
 1.79
3.46
12.6
Sb
Pb
100
35.6
14.5
W
Pb
100
1.59
3.4
Zn
Pb
250
3.18
0.11
Very large thermoelectric e.m.f.'s are observed in circuits composed
of semiconductors (Table 56).
TABLE 56
THERMOELECTRIC E.M.F. IN SEMICONDUCTORS
Range in C
(1)
(2)
a
b
From
To
Bi 2 0,
Pb
500
800
1946
186
Co,0 4
Pb
200
1200
629
 22.1
Cr 2 0a
Pb
950
1285
 704
43.2
CuO
Pb
170
850
 1029
171
PbO
Pb
250
390
48300
11720
Si
Pb
200
350
 408
 47
ZnO
Pb
355
1350
 735
25.$
As to the Thomson coefficients, their temperature dependence can
be represented by the empirical formula
<r  [a + 10 2 # + 10V] X 10* volt deg 1 . (20.55)
366
TEXTBOOK OF THERMODYNAMICS
XX 139
It is apparent from Table 57 that the Thomson effect is positive
in some metals and negative in others.
TABLE 57
THOMSON EFFECT
Range in C
Metal
a
ft
7
From
To
Ag
123
127
1.17
0.50
Al
 13
119
0.04
+0.475
Au
100
103
1.49
0.44
Bi
+ 25
32
+6.76
+2.8
Cu
 60
127
1.42
0.74
Fe
 51
115
44.00
+8.4
Pb
153
117
+0.61
+0.221
0.38
Pt
 72
128
+9.10
0.475
+4.75
Zn
173
40
2.74
1.15
Only the order of magnitude of the constants in Tables 55, 56, and
57 is significant: the thermoelectric power is highly sensitive to the
purity and mechanical treatment of the substances and shows con
siderable variation in different samples of the same material. Never
theless, it is possible to test the thermodynamical relation (20.50) by
carrying out on the same sample the measurement of the e.m.f. and
the calorimetric determinations of II and a. The agreement is very
good, as appears from the examples in Table 58.
TABLE 58
TEST OF THERMODYNAMICAL RELATIONS
n
dE
(T2 <TI
d*E
Metals
T
dT
Reference
T
dT 2
Reference
In 10
" fl volt
In 10
* volt
Cu Pt ...
3 66
3.67
(1)
3
3 6
(2)
Cu Fe
10.16
10.15
(1)
2
2 7
(2)
Cu nickeline
18.90
18.88
(i)
6.01
6.28
(2)
Cu German silver .
25.25
25.22
(i)
3.92
3.79
(2)
Borelius, Ann. Physik 56, p. 388, 1918.
* Berg, Ann. Physik 32, p. 477, 1910.
XX 139 MAGNETIC AND ELECTRIC PHENOMENA 367
Recent measurements of the Thomson coefficient at very low tem
temperatures l showed that <r vanishes in supraconductors and con
forms to eq. (20.53) in nonsupraconductive metals.
Let us now find the connection between the Peltier coefficient Hi2
(or the thermoelectric power Ui2/T) and the difference of potentials
12 1 122 which we calculated in sections 109 and 112 from the electron
theory of thermoelectricity. We must bear in mind, however, that in
its [present form this theory cannot yet
account for the detailed phenomena and
gives correctly only the order of magnitude
of the thermoelectric power. The results
of Chapter XVII refer to open chains, but,
if we take the case of a closed circuit (Fig.
62) with a very large resistance, the current
will be very weak and potentials com
pensating the thermoelectric forces will be FIG. 62. Thermoelectric
built up, practically, to the same extent as couple,
in the open circuit. Starting from eq.
(20.46), we can apply to the heat function the expressions it has in
equilibrium. We begin with the case when the electron gas is non
degenerate obeying eqs. (17.01) and (17.02) of the classical perfect
gas. The heat function has then the expression x = c p T + Wo + /ft,
so that we obtain, taking into consideration (17.13),
H 12 = 12 2 fli = log (20.56)
e z\
It was mentioned in section 109 that this expression applies only to
semiconductors. As k/e = 86 X 10 ~ 6 volt deg~ ! , this would imply
that log (z 2 /zi) for some of the substances of the Table 56 is of the
order of magnitude 20, whence z 2 /zi would be about 5 X 10 8 . While
it is within reason that a semiconductor should have an electronic
density 10 8 times smaller than that of lead, the signs of a in Table 56
offer a greater difficulty. They seem to indicate that in some of the
semiconductors the carriers of the electric current have a positive
charge. As we shall see in the next section this difficulty is not quite
so baffling as it was a few years ago.
Turning to the case of the degenerate electron gas, we shall make
use of the equation ^2 = ^1, from which the results of the section 112
were obtained. In view of the relation v = x Ts, we can write
1 Borelius, Keesom, and Johansson, Comm. Leiden, 196a; Borelius, Keesom,
Johansson and Linde, ibidem, 2l7a, b, c.
368 TEXTBOOK OF THERMODYNAMICS XX 140
/Hi2 X2 ~ xi = T($2 si). Hence with the help of eq. (17.33) we
obtain the formula due to Sommerfeld
_
1112 " 2
^ e ) Lft,2 fltiJ '
or
~ = 3.67 X 10~ 8 T\ ^  ^ ) volt deg 1 . (20.57)
1 \\li2 A*l/
For the combination leadsilver the formula (with the data of
Table 55) gives U 12 /T = 0.86 X 10~ 8 volt deg 1 , which is of the right
order of magnitude. However, the electron theory, in its present
form, is incomplete as it does not account for the phenomena quanti
tatively.
According to eq. (20.50), the e.m.f. is obtained from the thermo
electric power by the relation
E
/T
~f '
Intimately connected with thermoelectricity are the socalled
galvanomagnetic and thermomagnetic effects describing the influence of
a magnetic field on the phenomenon of electric conduction. The
scope of this book does not permit us to enter into their discussion,
and we refer the reader interested in the subject to its analysis in a
paper by Professor Bridgman. 1
140. Semiconductors. The difference between metallic con
ductors and semiconductors is not only quantitative but also qualita
tive. The conductivity of metals (in the range about C) has numeri
cal values between 10 15 and 10 17 <abs es and is inversely proportional
to the absolute temperature
Xccl/r,
so that it has a negative temperature coefficient. On the other hand,
in semiconductors \ is only of the order 10 4 to 10 13 and its temperature
dependence can be best represented by the law 2
X = Xoexp(a/r) (20.58)
with a positive temperature coefficient.
Recent work on the quantum theory of electrons in metals 3 has
1 P. W. Bridgman, Phys. Rev. 24, p. 644, 1924.
1 W. Voigt, Krystallphysik, 1910; E. Engelhard, Ann. Physik 17, p. 501, 1933.
* A. H. Wilson, Proc. Roy. Soc, 133, p. 458; 134, p. 277, 1931.
XX 140 MAGNETIC AND ELECTRIC PHENOMENA 369
greatly contributed to our understanding of the peculiarities of semi
conductors. As a part of the argumentation can be given a thermo
dynamical guise, 1 we shall outline here its main ideas, even though
the experimental data for testing the results are extremely meager.
While treating electron clouds in metals in Chapter VII, we
assumed that the electrons are free, actually implying in these words
two different assumptions. In the first place, the free electron does
not belong to any particular atom; it is shared by all atoms and can
travel from one end of the conductor to the other. In the second
place, we assumed that the energy levels of the free electrons can be
calculated, as if the atoms were absent and as if the electron gas were
contained in an empty box. It is of particular importance that the
energy levels in our calculation were distributed practically continu
ously: any small energy increment could raise a few of the electrons
to a higher level. We shall show now that electrons, which are free
in the first sense, need not necessarily be free in the second, and that
in this case they do not contribute to the conductivity, they are not
conduction electrons.
To fix our ideas let us consider two elements, the one divalent, the
other monovalent. The free divalent atom has two valency electrons
which, in their normal or 5states, have the same energy and differ
only by the orientation of their spins. There are no other states of
the same, or nearly the same, energy available: just as many 5states
as electrons. The next higher electron levels belong to the first
excited or state and have a considerably larger energy. It is differ
ent with the monovalent atom] its single valency electron can assume
a new quantum state without change of energy, simply reversing its
spin: there are twice as many 5states as electrons. When Z atoms
are brought together to form a crystal lattice, we know from general
mechanical principles the following two facts. (1) The number of
quantum states remains the same. The Z atoms had in their free
condition 2Z 5states; therefore, the number of 5states in the crystal
is also 2Z. (2) While the 5states of the free atoms had all the same
energy level, this level is split up in the crystal owing to the forces of
interaction between the atoms. In general it is replaced by 2Z
slightly different levels lying close together. In short, to the 5states
of the free atoms there corresponds in the crystal an 5band of closely
crowded levels (Fig. 63a).
The distribution of electrons over the states of the band is very
different in the cases of monovalent and divalent elements. In
J R. H. Fowler, Pfays. Zs. Sovjetunion 3, p. 507, 1933.
370 TEXTBOOK OF THERMODYNAMICS XX 140
monovalent elements there are only half as many electrons as quantum
states. Because of the exclusion principle, there can be only one
electron in each level so that, at T = 0, the electrons would fill the
lower half of the band. At finite temperatures, they can spread out
into the upper half. These are the conditions of freedom, in the
second sense mentioned above, which formed the basis of the electron
theory of section 111. It is immaterial that the band has an upper
limit because the electron gas is under ordinary conditions largely
degenerate and the overflow above the lower half is but small. On
the other hand, in divalent elements, there are in the sband exactly as
many levels as electrons. Supposing that the excited states (forming
(a)
FIG. 63. Position of s and bands in (a) insulators, (6) metallic conductors.
the />band) lie so high that they are unattainable at ordinary tem
peratures, each state is occupied by an electron: the electrons are
restricted to their positions in the band, they are " bound electrons 1 '
in the second sense. 1
What bearing have these facts on the problem of electric conduc
tion? From the point of view of the quantum theory, the mechanism
by which an electromotive force produces a current is redistributing
the electrons over the quantum states. According to the above analysis,
such a redistribution is possible in the case of a monovalent element
because it possesses plenty of unoccupied levels of accessible energy.
A monovalent element is, therefore, always a conductor (provided its
electrons are "free in the first sense"). On the other hand, a divalent
crystal may be either an insulator or a conductor, even when its elec
trons are shared by the whole system. When the pband of excited
levels lies so high as to be inaccessible (Fig. 63a) the electrons com
pletely fill the sband, they are " bound electrons" and their redistri
bution is impossible: the crystal is then an insulator. On the con
1 Interchanges of electrons need not be considered since they do not lead to a
new state of the system.
XX 140
MAGNETIC AND ELECTRIC PHENOMENA
371
impurity level
trary, when the pband lies low and partially overlaps with the 5band
(Fig. 63&), it can take the overflow of electrons so that they can be
redistributed: the crystal is a metallic conductor.
In this way, the quantum theory gives us a simple explanation of
why some substances are metals, others dielectrics. It has, more
over, the great advantage that the several kinds of semiconductors
can be easily fitted into the picture.
(A) Intrinsic semiconductors. Suppose that the pband of excited
states lies neither very high nor so low as to overlap with the 5band,
but has an intermediate position (Fig. 64a). We call Ae the energy
difference between the lowest />levels and the highest 5levels. We
consider the case that, at T = 0,
the 5band is completely filled with
electrons (as in the preceding ex
ample of divalent elements). At
higher temperatures, there is the
possibility of an electron spontane
ously leaving the 5band and rising
to one of the ^levels. We assume
that Aeis small enough for the prob
ability of such a rise being appre
ciable at room temperature, so that
the number of electrons belonging
to the pband is z_ per unit volume
of the crystal. These Z electrons
are free carriers of conductivity because there are plenty of unoccupied
states in the vicinity of their energy levels. Because of their removal,
there appears in the 5band an equal number (z+ = 3__) of unoccupied
levels or "holes." It was pointed out by Peierls l that the formulas
of the quantum theory can be given the following interpretation.
The "holes" behave in every way as if they were free carriers similar to
the electrons but having the opposite (positive) charge. In this way
we arrive at the following model of an intrinsic semiconductor: it
contains two clouds of conductive carriers, the cloud of free negative
electrons (G_) and the cloud of "free (positive) holes" (G+). The
"bound electrons" (G B ) of the 5band can be left out of consideration
as far as conduction phenomena are concerned.
A free electron of the pband may, of course, drop into a vacancy
("hole") of the 5band and so become a bound electron. This process
can go also in the opposite direction and can be described in the sym
1 R. Peierls, Zs. Physik S3, p. 255, 1929. Compare also: Bronstein, Phys. Zs.
Sovjetunion 2, p. 28, 1932.
(a) (b)
FIG. 64. Position of 5 and />bands
in (a) intrinsic, (b) impurity semi
conductors.
372 TEXTBOOK OF THERMODYNAMICS XX 140
bols of a chemical reaction as GL + G+ GB = 0. The equation of
equilibrium is, therefore, according to (17.09),
^ + (p+ <pB = 0.
The concentrations of free electrons and of free holes are very small
so that both clouds can be regarded as nondegenerate in the sense of
Chapter XVI. We shall assume, as a first approximation, that they
obey the equations of classical perfect gases. On the other hand, the
bound electrons are completely degenerate, and their thermodynamic
potential is reduced to a constant (section 111), namely the mean
energy level from which the transitions take place. If we measure
all energies from this level we can ignore <f> B altogether and have the
same conditions as in the case of equilibrium of positive and negative
electrons. Therefore, we can apply eqs. (19.28) and (19.29) of section
128 with the only difference that here the heat of reaction is Ae l
*_= 4.9 X 10 15 r % exp (11 600Ae/2r) (20.59)
(if Ae is expressed in electronvolts), a reasonable result as we shall
see below. 2 Since the electrons and "holes" are treated as classical
perfect gases, the electric conductivity can be computed by means of
the DrudeLorentz formula which was derived precisely under this
assumption : X = %e 2 (2irM e kT) " ^lz. The mean free path / is known to
be inversely proportional to the temperature, / = feo293/r (feo being
the value at 20 C). Substituting the numerical values of M e and jfe,
we can write, therefore,
X= 1.01 X10 6 r*/2os. (20.60)
Using the expression (20.59), and seeing that there are two kinds
of carriers, we obtain
X  1.0 X 10 21 /2oexp(5800Ae/r). (20.61)
As /2o is of the order 10 ~ 6 cm, the expression is of a reasonable
order of magnitude when Ae is of the order 0.1 ev.
(B) Impurity semiconductors. There exist substances which are
insulators in their pure states but semiconductors when impure.
Often as small an admixture of foreign atoms as 1 in 10 6 , or even less,
1 Already J. Koenigsberger (Jahrb. Radioaktiv. 4, p. 158, 1907) suggested an
exponential dependence of 2 on 1/3T.
2 A. H. Wilson (loc. cit.) does not treat the electrons as a perfect gas. His
formula for * is a little more general, differing from ours by a factor k/ft, where ft
measures the tightness of binding of the electrons and is a characteristic constant of
the material.
XX 140 MAGNETIC AND ELECTRIC PHENOMENA 373
is sufficient to produce an appreciable conductivity. Cuprous oxide
(Cu2O), the only semiconductor investigated with modern accuracy,
belongs to this class. The impurity on which the conduction of Cu2O
depends is free oxygen: it has been shown 1 that the conductivity of
this material is lowered by glowing in vacuo and increased by heating
in an oxygen atmosphere. The interpretation in terms of our model
is as follows: conduction electrons in the pband are supplied by the
impurity atoms. The distance between the p and $bands is so large
that in the pure substance the pband would be empty. On the other
hand, each impurity atom possesses an electron which can be more
easily separated because its energy level is only little (Ae) below the
lowest levels of the pband (Fig. 646). We denote by 0 the number
of conduction electrons in the pband, by z+ the number of ionized,
and by z of neutral, impurity atoms. In analogy with the preceding
case, the equation of reaction can be written GL + G<+ d = 0,
and the equation of equilibrium ^? + <?>+ ^< = 0. We regard
again the conduction electrons as a classical perfect gas and use for it
the expression (19.04). As to (pi = u Ts + pv of the impurity
atoms, they must be considered as stuck in the crystal lattice and
having a negligible thermokinetic motion: 2 therefore, u = o, po^ 0,
while their only entropy is the entropy of mixing given by eqs. (15.03)
of section 98 or (1 1 .39) of section 80, so that s t + $< = R log (z+i/zt).
Substituting the numerical values of the chemical constant and of Jfe,
logio(**<+/*<)  15.69 + f logio r  5040Ae/r,
or
* * 7.00 X lOVSr* exp( 5800Ae/jT). (20.62)
Vogt 3 and Engelhardt measured the coefficient of the exponential
(i.e. 7.00 X 10 7 \/ZiT^ in our theory) and obtained values between
1.5 X 10 17 and 9.4 X 10 19 in different samples. This would give for
Zi numbers from 1 X 10 16 to 4 X 10 20 , or from 4 X 10~ 7 to 2 X 10~ 2
impurity atoms per molecule of Cu2O. The measured Ae was about
0.65 ev. The conductivity, as obtained from (20.60), is
X  7.0 X 10 12 Zi"r*l2Q exp ( SSOOAe/T). (20.63)
(C) Lightsensitive semiconductors. There exist substances whose
conductivity increases many times when they are exposed to light,
1 R. Engelhardt, he. cit.
* A more rigorous statement is that the impurity atoms have the same thermo
kinetic motion whether they are ionized or not, so that the differences in u and pv
drop put of the equation.
W. Vogt., Ann. Physik 7, p. 183, 1930.
374 TEXTBOOK OF THERMODYNAMICS XX 140
the bestknown example being selenium. This case also fits readily
into the model of Fig. 63a. Here the pband lies so high that the
number of its electrons is negligible in the dark. However, when
light of sufficiently high frequency is absorbed by the crystal, it raises
electrons into the band by photoelectric action. The return into
the sband requires a radiative transition and may be a very slow
process, thus assuring an appreciable density of conduction electrons.
One of the most important aspects of the theory outlined above is
that it provides a new type of positive carriers of electric conduction :
the ''free holes" of the 5band. We have considered under (B) impuri
ties which supply free electrons. There may exist other impurities
which have an affinity for electrons and are able to bind them, thus
producing "holes". In this way some conductors may have an excess
of negative, others of positive, carriers. More important still, in the
more accurate theory the electrons and "holes" are not regarded as
entirely free. The binding of the two kinds of carriers is different,
and, for this reason, there is an asymmetry in their actions, even if
they are present in equal numbers. It was always emphasized by
E. H. Hall that the peculiarities of the effect bearing his name could
not be explained without an assumption of this sort. The same
hypothesis would be helpful in accounting for the great variations in
the sign of the thermoelectric e.m.f. and of the Thomson coefficient.
CHAPTER XXI
THE DIRECTION OF THERMODYNAMICAL PROCESSES
141. General remarks. General rules predicting the direction in
which a process is influenced by outer forces are very desirable from a
practical point of view. They may offer a quick orientation in the
workings of new experimental arrangements and so facilitate their
theoretical and practical understanding. This applies to thermo
dynamics even more than to other parts of physical science because
the thermodynamical treatment is more abstract and formal and less
accessible to visualization. The attempts to establish such rules,
with respect to thermodynamical processes, were all influenced,
directly or indirectly, by the famous principle of electrodynamics enun
ciated in 1833 by Lenz l : "When a force acting on a primary electric
current induces a secondary current, the direction of this secondary
current is such that its electrodynamical action opposes the acting
force."
An analogous principle of thermodynamics was formulated by
LeCha teller 2 in 1884 and in an extended form by Braun 3 in 1887.
Their procedure was essentially inductive (although Braun included a
few formulas in his paper) : they reviewed a large number of examples
which they tried to state in a form roughly analogous to Lenz's princi
ple. They claimed that all these examples could be regarded as special
cases of a general rule which they proceeded to formulate. However,
this formulation was so vague that it was impossible to apply the rule
without ambiguity. This fact was first pointed out by Raveau 4 and
Ehrenfest 5 and has been generally accepted since. It will serve no
purpose to mention here the form they gave to their principle, but it is
well to reproduce two of the examples from which it was derived.
1 H. F. Lenz, St. Petersburg Acad. of Sci. 29, XI, 1933; reprinted: Ann. Phys. u.
Chem. 31, p, 483, 1934.
2 H. LeChatelier, Comptes Rendus 99, p. 788, 1884.
1 F. Bra'un, Zs. phys. Chem. 1, p. 269, 1887; Ann. Physik 33, p. 337, 1888.
* M. C. Raveau, J. Phys. 8, p. 572, 1909.
' P. Ehrenfest, J. Russ. Phys. Soc. 41, p. 347, 1909; Zs. phys. Chemie, 77,
p. 227, 1911.
875
376 TEXTBOOK OF THERMODYNAMICS XXI 142
(A) Influence of pressure on solubility. Substances whose solubility
increases with pressure dissolve (at constant temperature) with volume
contraction; those which have the opposite pressure dependence dis
solve with volume dilatation. This fact (although not new) was
rediscovered by Braun a little earlier and was the starting point of his
considerations.
(B) Influence of temperature on compressibility. Suppose that we
compress a gas, increasing the pressure under which it stands by the
small amount A. We can do this in two different ways: (1) keeping
the temperature constant (T = const) with the help of a heat bath;
(2) "leaving the gas to itself" (i.e. adiabatically, 5 = const). The
compressibility is smaller in the second case
dp
dp
(21.01)
We selected the two examples so as to show the heterogeneity of
the cases considered by LeChatelier and Braun. The first involves
only three parameters of the system: pressure, volume, and concentra
tion of the saturated solution. In the second we have an interplay
of four parameters: pressure, volume, temperature and entropy.
The closer analysis will show that the two cases are also quite different
in their mathematical characteristics. In fact, Braun himself dis
cussed the two types represented by them separately but claimed, in
the end, that they are reducible to the same principle. Without
maintaining that all possible examples belong to the types (A) or (B),
we assert that the principle of LeChatelierBraun must be resolved into
at least two different and unconnected rules which we shall analyze in the
next two sections.
142. Displacement of equilibrium involving transformation of
matter. It will be noticed that example (A) of the preceding section
is only a special case of the first of the two laws stated at the end of
section 59. These laws, in their turn, are consequences of eqs. (6.51)
and (6.53) derived from the fundamental conditions of equilibrium,
as stated in section 40, namely:
dU = 0, (21.02)
with the subsidiary conditions
55 =0, 5F = 0. (21.03)
The gist of the argument which led in Chapter VI to the explicit
equilibrium conditions is as follows. The variables S and V possess
XXI 142 DIRECTION OF THERMODYNAMICAL PROCESSES 377
the additive property. In a heterogeneous system the total entropy
and the total volume are sums of the contributions of the several phases
S = S ( , V = 2] F<. (21 .04)
i J
As was shown in section 40, this property together with the condi
tions (21.02) and (21.03) immediately leads to the uniformity of the
parameters T and />, conjugate to S and V
(21.05)
The generalization for systems depending on other variables than 5
and V is obvious. For instance, in Chapter XX we discussed mag
netization and electrostatic polarization. We considered systems
placed in homogeneous magnetic or electric fields of the strength H
or E (i.e. being homogeneous and having this strength when the sys
tem is not there). The element of work was denoted by HdM 9
EdP, where M and P are the magnetic and electric moments of the
system in the direction of the fields. When the system consists of
several phases, the equations analogous to (21.05)
> = E,
are a matter of course since all the phases are placed in the same
homogeneous field. Moreover, the moments have the additive prop
erty
M = Z) Jf ( >\ P =><'>.
i J
Finally, it is known from the electromagnetic theory that in equi
librium
5M = 0, 6P 0.
The set of equations governing the properties of these variables is
formally identical with that applying to the parameters 5, V.
Passing to the general case we assume a heterogeneous system
which has the following properties:
(1) Each phase (j) is described by the variables
which have the additive property
EXi 01  JT* (21.06)
378 TEXTBOOK OF THERMODYNAMICS XXI 142
In particular let Xo (;) stand for the entropy 5 (j) .
(2) The differential of the internal energy has the form
y{fldX{*>. (21.07)
We call the coefficient yi (j) the generalized force conjugate to the
additive variable Xi ( >\ The temperature T (;) = y$ (j) is classed as a
generalized force j 1 the pressure p must be counted as a negative force
(p =  yi (i)).
(3) The equilibrium of the system is determined by
8U = 0,
with the subsidiary conditions
Mi** 0, (/ = 0, 1,. .. w). (21.08)
(4) The chemical composition and mass of each phase are deter
mined by the mol numbers Nk { ' } of its components (k = 1, 2, . . . 0).
The variations dN k (i) are given by the reactions possible in the system
(section 42): dN k <i} oc ^ (y) ,
S^C^G* ^ 0. (21.09)
The explicit equilibrium conditions can be obtained exactly in the
same way as in section 40 where a less general case was treated. The
procedure need not be repeated here but only indicated in its bare out
line. The problem is broken up in two partial ones: first, changes of
composition and mass are excluded (6JV* (y) = 0). The variation of the
internal energy has then the form
C 'W>=0, (21.10)
j i
with the subsidiary conditions (21.08). This gives immediately (as
in section 40)
yz (y) =yz, (/ = 0,1, . . .n). (21.11)
The generalized forces are uniform throughout the system. To treat
the second partial problem, we introduce the generalized thermo
dynamic potential
* (2112)
1 As compared with (2.02) the coefficients y are defined with the opposite sign.
XXI 142 DIRECTION OF THERMODYNAMICAL PROCESSES 379
whose differential is, according to (21.07),
d& =  *,< W>, (21.13)
i
so that it must be regarded as a function of the variables
We denote the partial thermodynamic potentials, as before,
When the generalized forces yi are kept constant, the conditions
(21.02) and (21.08) can be replaced by
a$ = o.
for the same reasons which were adduced in section 36.
We consider only variations in which yi = const; the condition
6$ = is then valid and can be written in the form
(; V ;) =0. (21.15)
j t
The partials of A< with respect to the generalized forces are, accord
ing to (21.13),
(21.16)
uyi
where
is the change of the variable Xi (for the whole system) in the trans
formation or reaction (21.09)
Apart from the generalized definition of 3> (/) the condition (21.15)
is precisely the same as (6.50) of section 42. The consequences derived
in Chapters VII to XI are based on this condition and remain true.
In particular, applying the generalized form of the thermodynamic
potential to the expressions (9.13) or (11.15) of the equilibrium con
stant, we find from (21.16)
9 log* **'
a* " RT (2L17)
The increase of the generalized force yi produces a displacement of
equilibrium attended by a process (transformation or chemical reaction)
in which the change of the conjugate variable AXi is positive.
380 TEXTBOOK OF THERMODYNAMICS XXI 143
In addition to example (A) of Braun's (preceding section), we shall
illustrate this principle by considering the influence of a magnetic field
on the equilibrium: yi = H, X t = M t
The increase of the magnetic strength of field produces a displace
ment of equilibrium in which the magnetic moment of the system (in
the direction of the field) is increased. The more paramagnetic phases
and components increase their mass at the expense of the less para
magnetic. Since supraconductors are classed as diamagnetics, this
includes also the equilibrium between metals in the normal and the
supraconductive states: a magnetic field makes the normal phase
increase at the expense of the supraconductive. These phenomena
are manifestations of the same causes which make an inhomogeneous
magnetic field suck in paramagnetic and push out diamagnetic bodies.
In a similar way, an electric field favors stronger dielectrics at the
expense of weaker ones.
143. The restricted LeChatelierBraun principle. (Influence of
secondary forces.) Example (B) of section 141 is a true analogue of
Lenz's principle: when the temperature is allowed to change, it opposes
the acting force of compression. This is in keeping with the expres
sion (21.07) in which the temperature is treated as a generalized force.
Another example of the same sort is the longitudinal compression of
of a solid elastic bar with uniform crosssection. It can be compressed
by a longitudinal force pi in two ways: (1) keeping the forces (of
pressure) on the lateral surfaces constant (p2 = const) ; (2) keeping the
lateral dimensions constant (/2 = const) and preventing a bulging out
of the crosssection. If the length of the bar is denoted by /i, it is
known from the theory of elasticity that
dpi
<
<2
dpi
(21.19)
Here, too, the secondary (lateral) force, if permitted to change,
opposes the acting (longitudinal) force. As it is sometimes said, the
secondary forces help the system to resist the acting force.
The analytical connection of the examples of this type with the
foundations of thermodynamics was cleared up by Ehrenfest. 1 We
shall call the rules pertaining to class (B) the restricted LeChatelier
1 See footnote on p. 375.
XXI 143 DIRECTION OF THERMODYNAMICAL PROCESSES 381
Braun principle. Ehrenfest himself refrained from formulating the
principle, 1 but the enunciation given below is based on his analysis.
Unlike our treatment in section 142, we may restrict our con
siderations to homogeneous systems described by the parameters
Xo, Xi, . . . X n so that the differential of the internal energy has the
expression
dU = y Q dX + yidXi + . . . + y n dX n , (21.20)
defining the generalized forces yo 9 yi, . . . y n  As dU is an exact dif
ferential, the following reciprocity relations hold:
In addition to them, we shall assume the validity of the inequalities
~ ^ 0, (21.22)
U** I
which are often called conditions of stability. If the variables X are
those of section 142, the first of these conditions is QT/QS ^ 0, which
expresses the fact that the temperature rises when heat is imparted to
the system. As to the condition relating to the volume (X\ = F), it
must be borne in mind that y\ = />, so that dp/QV 3* 0. In fact
the sign generally given to the hydrostatic pressure is an anomaly
among the generalized forces. In hydrodynamics and elasticity the
tension, and not the pressure, is taken as positive. The inequality
dp/dV ^ was discussed at length in section 5, where it was shown
to be a necessary condition of stability. The same applies in all cases
where the generalized force yi can be envisaged as different in the
system and in the environment. If the condition (21.22) were not
fulfilled, the difference would tend to increase and a stable equilibrium
between system and environment would be impossible. However, this
argument does not apply to cases like magnetization and polarization
(sections 134 and 138) where the generalized forces are defined as H
and E 9 as they would be in the absence of the system. Here the
validity of the conditions (21.22) is simply a coincidence, as far as
thermodynamics is concerned. As a matter of fact, diamagnetic bodies
form an exception to the rule: for them QH/QM < 0, if M is the mag
netization in the direction of the field. 2 Fortunately, the principle we
1 The reason for this was that he did not segregate class (A) into a separate
principle and was looking for a single rule embracing all cases. Planck (Ann. Physik
19, p. 759, 1934) also failed to draw the distinction.
* It would not help to define yi * M because this would reverse the sign of
i in eq. (21.20).
382 TEXTBOOK OF THERMODYNAMICS XXI 143
are going to formulate need not be amended because of diamagnetics
(see below). Sometimes it is said that the inequalities (21.22) hold
because the internal energy U has its minimum in equilibrium. But
such a conclusion may not be drawn when the minimum is contin
gent upon subsidiary conditions of the type (21.03). In this case,
the equations of stability (21.22) are not, in general, invariant with
respect to transformations of coordinates.
j All we shall need for the proof of the principle are the relations
(21.20), (21.21), and (21.22). It is immaterial whether the parameters
Xi have the additive property or not; therefore, we drop the requirement
of additivity. We are now ready for the enunciation.
(I) Suppose that the system is described by the variables Xo, . . . X n
and the generalized forces yo, . . . y n for which eqs. (21.20) and (21.21)
and the inequalities (21.22) are satisfied.
(II) We keep all the variables X constant except two, Xi and Xi, and
we give a primary increment dyi to the force yi which acts directly on its
conjugate variable Xi. We let this increment change Xi in two different
ways:
(a) keeping the (secondary) force y^ constant (yi = const) ;
(6) keeping the variable Xi constant (Xi = const) and allowing
the force yi to change.
(III) We assert that the primary increment 8yi is less effective in
case (b)
(21 . 23)
In other words t the secondary force yi, if permitted to act, opposes the
primary force yi.
The proof is simple: the generalized forces yi and y^ as all the
coefficients of the expression (21.20), are functions of the variables X.
But since only Xi and Xi are allowed to change, we need to bring in
evidence only these two parameters
y,  yi(X lt X<), yi  y(jr, f Xi). (21.24)
We can resolve the second equation with respect to X <
XiXi(Xi, yi ). (21.25)
Now dyi/dXi is the reciprocal of QXi/dVit so that the condition
(21.23) can be also written
XXI 144 DIRECTION OF THERMODYNAMICAL PROCESSES 383
According to the rules of partial differentiation
dyi
i.e. yi in (21.24) must be differentiated in so far as it contains
explicitly and in so far as it contains it through the medium of
From the second eq. (21.24) we find, as in section 3,
We substitute this, taking into account (21.21)
. (2127)
V ' '
This finishes the proof of the inequalities (21.26) and (21.23) since
all three partials (dyi/dXi) v ., (dyi/dXi) Xi , (dXi/dyi) Xt are positive
because of (21.22).
It will be seen that diamagnetic bodies do not contradict the
principle although they do not satisfy the conditions (21.22). In
fact, for them Xi = M, yi = II and (dM/QXi) = 0, since the mag
netic moment depends only on //, as was pointed out in section 134.
For diamagnetics the second term on the right side of (21.27) vanishes
and they fulfill the limiting case of (21.23) corresponding to the sign
of equality. It is conceivable, of course, that applications may arise
in which the stability conditions (21.23) are not satisfied any more
than in diamagnetics but which are not so innocuous and break
through the law (21.23). It is obvious how the principle can be
extended to take care of such contingencies: one must write the
inequality (21.23) for the absolute values of the partials and must
make the sense (< or >) depend on whether (dyi/QXi) and (Qyt/QXi)
have the same or opposite signs.
144. Intensive and extensive quantities. The principle of
LeChatelierBraun is sometimes brought in connection with the
division of physical variables into " Quantitaeten " and " Inten
sitaeten " proposed by the defunct school of energeticists. 1 This
nomenclature is still used by German and Dutch writers and has
some relationship to the English terms: extensive and intensive quan
tities. As the reader is likely to come across both designations in his
l Main proponents: Mach, Ostwald, Helm. Compare G. Helm, Energetik.
Leipzig 1898.
384 TEXTBOOK OF THERMODYNAMICS XXI 144
study of foreign and domestic literature, it will be well to explain here,
briefly, their similarities and dissimilarities.
As defined by Tolman, 1 extensive quantities are those which have
the additive property, for instance, the variables Xo, . . . X n of section
142. Every quantity which is not additive is called by him intensive.
The division is intended to be exhaustive so that every variable must
fall into one of these two classes. It is possible, and often convenient,
to use as parameters describing the system not the variables X but
simple functions of them (like the reciprocals 1/X or the squares X 2 )
which are not additive. In this case, the system would be described
only by intensive variables. It must be said, however, that Tolman's
definitions are not universally accepted and that some authors use the
term intensive in the same sense in which we have used the words
specific quantity in section 39 (i.e. a homogeneous function of the
degree zero in the mol numbers).
On the other hand, the classification into Quantitaeten and Inten
sitaeten is not meant to be exhaustive. They are exceptional or
preferred variables, and a parameter chosen at random need not
belong to either of the two groups. Ostwald and Helm give no sharp
definition, saying only that the element of work (section 7) can be
represented as a sum of terms, each the product of a " Quantitaet "
and an "Intensitaet". It is safe to say that the additive variables X
of section 142 (i.e. the extensive quantities) belong to the " Quanti
taeten " and the generalized forces y, conjugate to them, to the
" Intensitaeten". But it is vague what other parameters (if any)
may be classed in these groups. The only practicable suggestion seems
to be that of Mesdames EhrenfestAfanassiewa and De HaasLorentz, 2
who propose to restrict these terms to such variables and conjugate
generalized forces which leave invariant eqs. (21.20) and (21.22) of
the preceding section.
1 R. C. Tolman, Phys. Rev. 9, p. 234, 1917.
1 T. EhrenfestAfanassiewa and G. L. DeHaasLorentz, Physica 2, p. 743, 1935.
CHAPTER XXII
LIMITATIONS OF THERMODYNAMICS 1
145. The statistical point of view. The statistical interpretation
of the entropy concept was treated in section 30. It was stated there
that the entropy of a thermodynamical system can be brought in con
nection with its probability P by means of Boltzmann's principle (4.67)
5 = klogP. (22.01)
While we emphasized in section 30 the parallelism between the
thermodynamical and the statistical points of view, we shall dwell now
on the discrepancies between them which have produced in the last
decades a profound change in the outlook of science upon the theory of
heat and, especially, upon the second law.
The identification of entropy and probability involves the following
difficulty. According to the second law, the entropy of an adiabatically
isolated system can never decrease: AS ^ 0. If we leave such a
system to itself, its entropy will either increase monotonically or
remain stationary. On the other hand, the probability P shows a
different behavior. Owing to the interplay of atomic and molecular
movements and forces, the system undergoes, continuously, small but
erratic changes. In general they occur in the direction of more probable
states. However, this is only an average effect: it lies in the nature of
the concept of probability that the less probable states are not totally
excluded but only less frequent in their occurrence. Therefore, the
probability of a system will, occasionally, decrease. In fact, it was
rigorously proved by Poincar 2 that a finite system, subject to the
Hamiltonian equations of dynamics, returns again and again to any
state through which it once has passed.
When Boltzmann's principle as expressed in eq. (22.01) was first
advanced, scientific thought divided itself into two schools. One school
1 With the kind permission of Yale University Press parts of this chapter were
patterned after the exposition by P. S. Epstein, Commentary on the Scientific Writ
ings of J. W. Gibbs, Article 0, Section 16.
* H. Poincare, Acta Mathematica 13, p. 1, 1890.
385
386 TEXTBOOK OF THERMODYNAMICS XXII 145
regarded the lack of conformity between thermodynamics and the
kinetic theory as a serious objection against the statistical interpreta
tion. The other contended that the principle of the increase of entropy
had itself only a statistical validity and was true only in the time
average. As Gibbs put it as early as 1876: "The impossibility
of an uncompensated decrease of the entropy seems reduced to an
improbability." The views of the two schools were brought to a focus
in a very interesting polemic between Zermelo 1 and Boltzmann. 2
The wide attention which this discussion received stimulated new work
on the subject, and, before long, the controversy was definitely settled
in favor of the statistical point of view through the investigations of
Von Smoluchowski 3 and of Einstein. 4 These authors showed that
deviations from the entropy principle do, actually, occur and can be
observed provided the system is sufficiently small. They accomplished
this by turning eq. (22.01) into a heuristic method for the treatment of
problems lying outside the scope of classical thermodynamics. Let /
be one of the parameters describing the state of the system, and let /o
be its value in the normal state of maximum entropy 5. We ask the
following question: what is the probability of this parameter assum
ing a value between /o + A/ and /o + A/ + dl? Let the entropy,
corresponding to /o + A/, be So + AS, where AS is, necessarily, nega
tive, So being the possible maximum. Inverting the relation (22.01)
between entropy and probability, we find that P must be proportional
to exp [(So + AS)/]. On the other hand, the probability must be
also proportional to the interval dl. Therefore, we can write for it
Pidl = Ce* s/ *dl. (22.02)
The factor C may depend on the normal values of all the parame
ters of the system. 6 It is determined by the condition
Pidl  1, (22.03)
which expresses the fact that the value of / will lie, certainly, in one
of the intervals dl.
E. Zermelo, Ann. Physik 57, p. 485, 1896; 59, p. 743, 1897.
1 L. Boltzmann, Ann. Physik 57, p. 773, 1896; 60, p. 392, 1897.
1 M. Von Smoluchowski, BoltzmannFestschrift, p. 626. Leipzig 1904; Ann.
Physik 21, p. 756, 1906; 25, p. 205, 1908.
4 A. Einstein, Ann. Physik 17, p. 549, 1905; 19, p. 373, 1906; 33, p. 1275, 1910.
* Strictly speaking, the parameter / must be selected in a definite way to make C
depend only on the normal values /o and not on the / themselves. We cannot enter
here into the rules of selection as they represent a problem of statistical mechanics.
XXII 146 LIMITATIONS OF THERMODYNAMICS 387
146. Brownian movements. Classical thermodynamics cannot
account for the fact that small particles dispersed in a liquid or gas
form a permanent suspension, that is to say, float in the medium and
are in a state of irregular agitation called Brownian movement. From
the point of view of that theory the particle is merely a part of the
boundary of the liquid, and its position is determined not by thermo
dynamical but by dynamical laws. These laws tell us that the position
of equilibrium for the particle is its motionless state at the bottom of
the vessel. Let us suppose the vessel to be of heatinsulating material
so that our thermodynamical system (liquid or gas) is adiabatically
isolated. In this case, the dynamical equilibrium of the particle
corresponds also to the maximum of entropy of the liquid (or gas). In
fact, it would require the energy expenditure e = M'gx to raise the
particle from the bottom to the height x, g being the acceleration of the
gravity field and M' the effective mass of the particle (its true mass
minus the mass of the displaced medium). This energy can be supplied
only by the liquid (or gas) which undergoes a corresponding loss of
internal energy A 7 = e. In general the entropy is expressed by
AS = (AZ7 + &W)/T\ in our particular case, the volume of the
liquid remains unchanged when the particle is raised and no work is
done by the system (AW = 0). Therefore, AS = e/T: this entropy
change is negative, and the strict point of view on the second law
would deny the possibility of the particle being permanently afloat.
On the other hand, the statistical point of view admits the existence
of small entropy fluctuations. The probability of the floating particle's
being found at a height between x and x + dx is given directly by the
formula (22.02)
. _M*g*
P x dx = Ce trdx = Ce *** dx. (22.04)
It is equally easy to find the probability for the particle being in
motion. Suppose the components of its velocity c lie between the
limits c z and c x + dc x , c v and c y + dc v , c g and c g + dc f while its kinetic
energy is Mc 2 /2. We conclude, as in the preceding case, that this
energy is supplied by the liquid which experiences the changes
AJ7 =  Mc 2 /2, AS =  Mc*/2T. The formula (22.02) gives for the
probability of this state of the particle
Pdctdc^c,  C f exp (Mc 2 /2kT)dc x dc v dc t . (22.05)
If the suspension consists of a large number of identical particles,
each having the mass Af, the formulas (22.04) and (22.05) determine
their numbers z at the height x and in the velocity interval
388 TEXTBOOK OF THERMODYNAMICS XXII 146
The first equation can be written
z = zo exp (M'gx/kT), (22.06)
where z is the number of particles (referred to unit volume) at the
height x, and zo at the bottom (x = 0). It is identical with the baro
metric law of Laplace, which was deduced in section 107 in a generalized
form starting from the law of perfect gases. We see now that the
particles of a suspension also obey the barometric law, and this implies
that their osmotic pressure is that of perfect gases: pv = RT or
p =5 zkT (section 66). The simplest way of showing this is to invert
the usual form of the derivation of Laplace's formula. At the height x
the pressure is higher than at x + dx by the amount dp because of the
weight of the particles in the layer dx. This weight is dp = zM'gdx
per unit area. On the other hand, we find from (20.06) x = kT
X log (Z/ZQ)/ M'g and dx = kTdz/zM'g, whence p = zkT.
The formula (22.05) represents Maxwell's velocity distribution and
is equivalent to the caloric equation of state u = %RT + #o. This
follows immediately from calculating the total kinetic energy of all
particles by integrating \Mf f J 9 c 2 Pdc ai dCj / dc t and keeping in mind
thatffrpdcyidcydcg = 1. Of course, the particles have also rotational
kinetic energy which is not included in our calculation of u. It is
easy to show that its amount is RT per mol, but the above examples
will suffice.
Extensive experimental work on suspensions was done by Perrin l
and Svedberg. 2 These authors tested the barometric formula (20.06)
directly, and Maxwell's distribution law (20.05) indirectly, and found
both in excellent agreement with experiments.
Brownian movements affect not only particles in a suspension
but indeed every movable object. For instance, a delicate tor
sion balance continuously undergoes small deflections from its zero
position. If the directing force is a and the angle of deflection tf,
the potential energy of the balance becomes e tf = ^ad 2 . As in the
case just treated, this energy is taken from the surrounding medium
(air) whose internal energy and entropy experience the changes
A 7 =  atf 2 , AS =  Ja# 2 /T. Again we find from (20.02) as the
probability of a deflection between tf and & + d&
Pjto = Cexp (%atf*/kr) d&. (20.07)
1 J. Perrin, Brownian Movement and Molecular Reality. London 1910.
* The Svedberg, Die Existenz der Molekflle. Leipzig 1912.
XXII 147 LIMITATIONS OF THERMODYNAMICS 389
The mean potential energy of the balance is, therefore,
/+
EtPadd.
J>
Determining the coefficient from the condition (22.03) we find
2e  ^ kT, (20.08)
corresponding to equipartition of energy (section 30). Observation
on a torsion balance were carried out by Kappler. 1 He found that
the distribution of the deflections is accurately represented by the
Gaussian curve (20.07). Substituting into (20.08) the measured mean
at? 2 , he obtained an experimental value for k and derived from it
(because of n A = R/k) the result n A = (6.06 0.06) X 10 ~ 23 mol" 1
for the Avogadro number.
147, Theory of fluctuations. According to classical thermody
namics, any substance is in equilibrium quite uniform and not subject
to spontaneous changes. The statistical view on the second law, on
the contrary, permits us to put the question as to local deviations
from uniformity. Let us focus our attention on a small part of the
system, a group of adjacent molecules, and let us calculate the proba
bility that the parameter / has in it an abnormal value. We denote
by MI, M2 the masses of the small part and of the remaining large
part of the system (M 2 MI), by so, k the normal specific values
of the entropy and of the parameter /. We consider the case, however,
when these quantities have in the two parts the slightly abnormal
values so + Asi, h + A/i, and SQ + As 2 , / + A/ 2 . The total devia
tion of entropy is, therefore, AS = MiAsi + M 2 As 2 or expanding
into a Taylor series with respect to A/i, A/2 (which we assume to be
very small quantities)
AS  ^
3A)
Since the entropy has its maximum in the normal state, the first
term must vanish: AfiA/i + Af 2 A/ 2 =0 or A/2 = (Mi /M 2) A/I.
In the second term M 2 (A/ 2 ) 2 = Mi(A/i) 2 Afi/Af 2 becomes then
negligible, whence
(22.09)
The entropy deviation in the small part of the system is expressed
1 E. Kappler, Ann. Physik 11, p. 233, 1931; 15, p. 545, 1932.
390 TEXTBOOK OF THERMODYNAMICS XXII 148
in terms of its own properties. We may, therefore, drop the subscript
(1) treating it as if it were alone. Equation (22.02) takes the form
Pidl = Cexp
The mean quadratic deviation of the parameter / (in the small
part of the system) is by definition
/^fo
(S/) 5 = C I
*/_00
(A/) 2 P,<tf,
while C is to be determined from (22.03). It must be remembered
that Qs 2 /Ql 2 is necessarily negative: S being a maximum, the deviation
can have only the negative sign. Therefore, the integration is easily
carried out giving
(22  10)
148. Density fluctuations and light scattering. We shall apply
this formula to compute the magnitude of the spontaneous density
fluctuations in a gas. Let us take as the parameter the molal volume
in the small selected portion of the gas: / = v. Considering the
fluctuations of one parameter only implies that the others drop out
in the averaging process, being on the average constant. The ques
tion arises, therefore, which other parameter we have to regard as
constant in taking the partial of 5 with respect to v. The answer is
simple in this particular case because we know from section 31 that a
subsidiary condition of equilibrium is dU = 0. Unless we want to
calculate the fluctuations of the internal energy itself, we must always
consider changes at u = const. We have, therefore, for the relative
fluctuation
/A,\2 / A A2 I.
(22.11)
where p oc i/v is the density, whence Ap oc Av/t^. From (4.15),
(9s/9i/)t = P/T, the result of the second differentiation at u = const
is little different from that at T = 0. We may write with a good
approximation (d 2 s/dv*)u = (dp/dv) T /T, or denoting the compressi
bility _^^
^' (22.12)
XXII 148 LIMITATIONS OF THERMODYNAMICS 391
In the case of the ideal gas, /3 =!//>, V = kZT/p, Z being the
total number of molecules in the considered small portion of the gas.
We obtain, therefore, _
/A.\2 1
(22.13)
as found already by Bernoulli for the fluctuations of independent units
(molecules). This formula shows us that the mean relative deviation
of the density from its normal value is very small for a large mass of gas
but becomes appreciable when the mass decreases. Measuring the
density of 1 cm of air (Z = 2.7 X 10 19 ) we are not likely to find any
deviation from the norm, as it is in the mean only 2 X 10~ 10 . How
ever, the deviation is measurable in an airfilled cube whose edge is
of the order of a wave length of light (5 X 10~ 5 cm) because here
Z = 3.6 X 10 6 whence 1/Z W = 0.53 X 10~ 3 . There exist, therefore,
in the gas numerous condensed and rarefied regions of this size and
smaller. This inhomogeneity must produce a scattering action on
light passing through the gas, and the amount of scattering to be
expected on the basis of the formula (22.12) was computed by Von
Smoluchowski l and by Einstein. 2
A simple way of deriving this formula is as follows. 3 The theory
of optics gives the following expression for the intensity of light
scattered by a small dielectric particle under the angle # from the
direction of the incident plane wave (whose intensity is /) and at the
distance r from the particle
AJ, ^F2(AD)2 1 + cos* tf
~T ~ ~~ (22 ' 14)
(Rayleigh scattering). AZ> is the difference between the dielectric
constants of the particle and of the surrounding medium. It is sup
posed that AD is small ; under this assumption the scattered intensity
depends only on the volume V of the particle and not on its shape.
Ao is the wave length of the incident light, as it would be in vacua. The
formula applies only to particles whose dimensions are small compared
with the wave length. We can regard the spontaneous inhomogeneities
of a gas as particles of this sort. Its whole bulk consists of such
inhomogeneities lying side by side, but if we focus our attention on one
of them, we can say that the average dielectric constant of its environ
1 See footnote on p. 386.
2 See last paper in footnote on p. 386.
8 P. S. Epstein, EnzyklopSdie der math. Wiss. V 3, p. 519, Leipzig, 1915.
392 TEXTBOOK OF THERMODYNAMICS XXII 148
merit has the normal value D while its own dielectric constant deviates
from it by AD. This deviation is connected with the density fluctua
tion (22.12) by the wellknown formula (D  !)/(!> + 2)p const,
whence
^ = A, P 1)^ + 2). (22 15)
P 3
We now put the question as to the scattering of a mass of gas of
the volume V M . The inhomogeneities are distributed in it quite irregu
larly, and there are no phase relations between the amounts of light
scattered by them. The total scattering is, therefore, simply the sum
of these amounts. According to our previous analysis the inhomo
geneities of small volume greatly dominate, so that there is no objec
tion to applying the formula (22.14): The summation of the expres
sions for all the homogeneities in V M will lead to F 2 (AZ>) 2 being
replaced by
Substituting from (22.15) and (22.12)
Using the abbreviation
8JT 3
a = ^ kT ^ D ~ " 1 ) S ( Z? + 2 > 2 ' < 22 ' 16 >
the scattering by a unit volume of gas becomes
1  L c2 '\ (22.17)
(22.18)
"" 16* r 2
or integrated over all space directions
 a.
a has, therefore, the meaning of the coefficient of extinction through
scattering.
In the particular case of permanent gases D is very little different
from 1, ft is practically 1/p, and kTft = 1/z, where z is the number of
molecules per unit volume:
Sir 3
 (D  I) 2 . (22.19)
XXII 148 LIMITATIONS OF THERMODYNAMICS 393
The last three formulas were first given by Lord Rayleigh 1 in his
theory of atmospheric scattering. He explained the blue color of the
sky as a consequence of the factor 1/Ao 4 : light of short wave length is
scattered more strongly so that the sunlight which indirectly reaches
our eyes, after being scattered in the upper layers of the atmosphere,
has its intensity maximum in the blue. Numerous measurements
confirmed the dependence on the wave length and on the angle #
given by these formulas. 2 The direct observations of the coefficient of
extinction give Sit*(D  l) 2 /3z == 1.00 X 10 ~ 24 cm 1 , reduced to 1 atm
and C, the calculated value being 1.04 X 10 ~ 24 cm 1 . The more
general formula (22. 16) shows that light scattering must be particularly
strong when the gas is in the vicinity of the critical state because the
compressibility $ is then very large. This explains the phenomenon of
the socalled opalescence in the critical state. It has been investigated
quantitatively by Keesom 3 and found in agreement with the theory.
In conclusion, we repeat that the entropy principle cannot account
for Brownian movements or density fluctuations. The existence of
these phenomena is conclusive proof that an uncompensated decrease
of the entropy is not impossible but only highly improbable. This
fact does not impair the value of thermodynamics as a method of
analyzing physical reality: in systems of appreciable size the change
of a deviation from the second law is so extremely small as to be
entirely negligible.
*Lord Rayleigh, Phil. Mag. 41, pp. 107, 274, 447, 1871; 12, p. 81, 1881; 44,
p. 28, 1897; 47, p. 375, 1899.
2 H. Dember, Ber. Saxon Acad. 64, p. 289, 1912; F. E. Fowle, Astrophys. J. Obs.
Smithsonian Inst. Ill, Washington 1914; E. Kron, Ann. Physik 45, p. 377, 1914.
* W. H. Keesom, Ann. Physik 35, p. 591, 1911.
APPENDIX I
COMPARATIVE TABLE OF NOTATIONS
There are in use four major and numerous minor systems of thermodynamical
notations. The major systems are as follows: (1) the Gibbs notations; (2) the
system developed by the physicists Clausius, Helmholtz, and Planck (C.H.P.);
(3) the notations adopted by the larger part of the American chemists, especially
Lewis and Randall (L.R.); (4) the system used in technical thermodynamics.
We believe that this lack of unity is neither strange nor particularly regrettable.
Thermodynamics is primarily a method for deriving relations between measured
quantities. Various branches of science make use of it, each applies it to its own
distinctive material, and in each the problem of harmonizing the notations and
avoiding duplications and ambiguities is entirely different. It would be extremely
difficult, if not impossible, to devise a system satisfying all these complicated require
ments. On the other hand, the inconvenience to the reader caused by the lack of
unity is slight, as the relation of the different systems of notations can be set down
in a simple comparative table.
As to the intrinsic merits of the different systems there is not much to choose
between them. A systematic principle was first introduced in the C.H.P. system
in the form of a rule with respect to additive or extensive quantities (sections 39
and 144): the total quantities (i.e. those referred to the whole system or, at least, to
a phase) are denoted by capital, the molal and specific quantities by lowercase
letters. L.R. modify this by using for molal quantities a special size of type, an
expedient which seems less convenient in classroom teaching.
We seriously considered adopting the L.R. system, which has a large following
among the chemists of this country. But their choice of the notations E and H
for the internal energy and heat function (enthalpy) is too repugnant to the physicist
who has to treat electric and magnetic applications of thermodynamics and is
accustomed to denote by E and H the electromagnetic field strengths. Worse still,
according to the above rule the molal energy and enthalpy would have to be denoted
by e and h, letters which are already badly overtaxed. We have, therefore, chosen
to use the C.H.P. system with the following slight corrections. (1) The letter F
for the work function ( U TS) seemed undesirable as it is used by L.R. for the func
tion U TS f pV; therefore, we substituted for it *, the choice of Gibbs. (2) The
use of UK for the mol number (i.e. number of mols of the component h in a system)
and of No for Avogadro's number is inconsistent with the above rule since the first
quantity is total and the second molal. We write, therefore Nh and n A .
In the following table we list the more important notations, omitting those which
are identical in all systems.
395
396
APPENDIX
Quantity
E.
C.H.P.
Gibbs
L.R.
Total
Molal
Total
Molal
Internal energy
U
X
s
*
$
N h
u
X
s
t
<f>
?A
nth
n A
U
X
5
F
r*
nn
u
X
s
f
TV
7V/i
X
y
t
r
Mfc
nh
E
H
S
A
F
*h
n h
Wh
N
Heat function (enthalpy)
Entropy
Work function (UTS)
Thermodynamic potential
(UTS+pV)
Partial molal thermodynamic potential
Mol number
Molality
Avogadro's number
No
....
Mol fraction
Xh
Xh,
Ch
N A
APPENDIX II
CONSTANTS AND CONVERSION FACTORS 1
TABLE OF CONVERSION FACTORS
1 norm, atmosphere
1 013 249 X 10 G dyne cm" 2
logio
6 005 717
1 (15) gcalorie
4.1852 joule
621 720
t <
4 1852 X 10 7 erg
7 621 720
< i
426 78 g m
2 630 199
1 00 095 (20) gcalorie
OftOO 4.1?
4 1
99 976 (mean) gcalorie
T 999 896
1 ev. (electronvolt)
2 0017 X 10~ 12 erg
15 30 140
n A ev
1 2139 X 10 12 erg mol"" 1
1? OS 4.10
<
29 003 cal mol ~ l
4 46 245
1 Partially based on the critical compilations by R. T. Birge, Phys. Rev. Supple
ment 1, p. 1, 1929; Phys. Rev. 40, p. 207, 1932.
398
APPENDIX
9
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fe c
c
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OOOOOOOOOOOOOOOOOO
o
TH
X
ISIS
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r r o
^ ^ * MM
1 1 1 I. 1 1 5
ooooooooo
XXXX XXXXXXXXX
&
I
D
d.
2

i
a
!
.,<
g
r
i!
'
3J
2
SUBJECT INDEX
Absolute entropy, see: entropy constant
Absolute temperature, 8, 58, 74
realization of scale, 7476
Absorption of gases in liquids, 161, 162
coefficient of , 161
Activity, 199, 287
coefficient, 201, 293, 294
function of electrolytes, 203205,
290, 294, 295
of gas mixtures, 205208
Additive property of entropy, 59, 60
Additive quantities, 103, 377, 384
Adiabatic, equation of , 48
in gas mixtures, 320
in perfect gases, 49, 263
in simple systems, 320
Adiabatic process, 48, 49
Adsorption in capillary layer, 223
Affinity, chemical , 92, 93, 225
Air, composition of , 308
specific heat of , 309
Annihilation of matter, 336, 343, 344
Atmospheric scattering, 392, 393
Auxiliary constraints, 99102
Avogadro's number, 6, 7, 398
B
Barrett effect, 350
Beattie and Bridgeman, eq. of state
of , 16
Binary gas mixtures, 13, 205208
Binary systems, 179194
Black radiation, energy and entropy
of, 331333
Boiling point, 116123
elevation of due to presence of sol
utes, 166, 167, 202
influence of surface tension on ,
217, 218
Boiling pressure, 11, 116123
lowering of by solutes, 163, 166, 202
influence of capillary layer on ,
216,217
of water, 1 19
Boiling temperature, see: boiling point
Boltzmann's constant, 7, 78, 398
Boltzmann's principle, 78, 79, 260, 296,
385
Born and Stern, formula for surface
tension of solids of , 220
Brownian movements, 249, 387389.
Caloric eq. of state (see: eq. of state), 40
Caloric properties of matter, 27
Calorie, 26, 397
Capillarity, 209224
Capillary layer, 210, 213
Capillary rise, 215
Carbon dioxide, curve of fusion and
sublimation of , 127, 128
isothermals of , 17
Carnot's cyclic process, 5052, 5456
efficiency of , 52.
Centigrade scale of temperatures, 5
Cells, galvanic (see also: galvanic cells),
94, 95, 287, 288
Characteristic functions, 38, 46, 62, 85
construction of , 85
Charged gases, 266293
equilibrium in , 268270
Chemical affinity, 92, 93, 225
Chemical compound, 149
Chemical constant (see also: entropy
constant of gases and Sackur
Tetrode formula), 236238
of degenerate perfect gases, 257
of diatomic gases, 313314
of electron gas, 278
of metallic ions, 284
of monatomic gases, 312, 313
of polyatomic gases, 314
399
400
SUBJECT INDEX
Chemical equilibrium,
general conditions of , 112, 113
in charged gases, 270
in perfect gases, 137147, 239
in processes involving radiation,
335, 336
in terms of activities, 200, 201
influence of electric field on , 270
influence of gravitational field on ,
271, 344
of solutes in dilute solutions,
176177
Chlorine, specific heat of , 308, 309, 3 1 1
ClapeyronClausius equation, 118, 120
Compressibility, coefficient of , 3, 8
in critical state, 12, 393
Concentration cells, 288
Concentrations of ions, 288
Condensation of the EinsteinBose de
generate gas, 258, 259
Condensation, process of , 218
Conservation of energy, see: first law of
thermodynamics
Constant, of entropy, see: entropy
of internal energy, see: internal
energy
Constants, table of fundamental , 398
Contact potential, 275
classical theory of , 273286
quantum theory of , 283
Continuity of liquid and gaseous states,
9,11
Conversion factors (see also: efficiency),
52, 397
Corresponding states, 12
extended law of , 68, 69, 123125,
221
law of , 12, 16
Covolume, 9
Critical constants, 12, 14, 15
Critical data, 14, 15
Critical state, 12, 13, 127, 128
opalescence in , 393
Cryohydrates, 188
Crystals, entropy of mixed , 77, 78,
246, 248
D
Debye's formula for specific heat of
solids, 324
DebyeHuckel, theory of strong electro
lytes, 290295
limiting law of , 294
Degenerate electron gas, 277, 278, 368
Degenerate perfect gases, 252265
numerical amount of degeneration in
, 263265
Density fluctuations, 390392
Deuterium, specific heat of , 306, 307
equilibrium of and hydrogen, 340
Diamagnetism, 347, 351
Dielectrics, 359, 360
Differential, exact , 20
Differential expression, linear , 20, 21
Differential JouleThomson effect, 70
Dilute solutions, 152177
conditions of equilibrium in , 155
157
Displacement of equilibrium, 156, 376
380
Dissociation, degree of , 140
depression of by various agents,
144148, 178
Distillation, 183
Distribution law, 157, 158
Doublet heat, see: specific heat
Duhem's equation, 103
Dulong and Petit, law of , 45, 232
Dushman's formula for thermionic cur
rents, 279
Efficiency, of Carnot's process, 52
of galvanic cells, 95
of heat engines, 51, 52, 55, 95
EinsteinBose gas, see: degenerate per
fect gases
condensation of , 258, 259
Einstein's formula for specific heat of
solids, 233
Einstein's law of equivalence of mass
and energy, 330
Electric field, influence of on equilib
rium in charged gases, 270
influence of on equilibrium of drop
lets, 219
Electrolytes (see also: weak and strong
electrolytes, and activity function),
94, 204
specific heat of , 329
SUBJECT INDEX
401
Electromotive force, 94
of galvanic cells, 94, 289, 290
of thermoelectric couples, 364, 365,
368
Electron clouds, 266285
Electron gas, 266285
degenerate , 277, 278, 368
Electron pairs, 337, 338
Electronic specific heat in metals, 278,
326
Electrons, positive , see: positrons
equilibrium of positive and negative
, 337, 338
Electrostriction, 360
Elevation of boiling point by solutes,
166, 167, 202
Endothermic process, 27
Energy, 27
conservation of , see: first law of
thermodynamics
internal , see: internal energy
of black radiation, 333
Enthalpy, see: heat function
Entropy, 58, 59, 60
additivity of , 59, 60
measurement of , 66, 229, 230
of black radiation, 333
of degenerate perfect gas, 256,
257
of dilute and perfect solutions,
153, 154
of electron gas, 278
of gas mixtures, 136, 207
of mixed crystals, 77, 78, 246,
248
of perfect gases, 63
of supercooled liquids, 248
of Van der Waals gases, 65
statistical interpretation of , 7679,
385, 386
Entropy constant, 59, 230
of gases (see also: chemical con
stant), 235, 236, 337
Entropy principle, 60, 61, 82
Eotvos* formula for surface tension,
221
Equation of state (caloric), 40, 68, 347
of degenerate perfect gases, 255
of perfect gases, 43
of Van der Waals gases, 65
Equation of state (magnetic), 347
of LangevinBrillouin, 348
Equation of state (thermal), 3
of Beattie and Bridgeman, 16
of Kamerlingh Onnes, 16
of degenerate perfect gases, 252, 253
of perfect gases, 6, 7
of Van der Waals, 12
of Van der WaalsLorentz for
binary gas mixtures, 13, 206
Equilibrium, 1
chemical , see: chemical equilibrium
displacement of , 156, 376380
fundamental conditions of , 82, 92,
93, 376379
influence of capillary forces on ,
213218
influence of electric forces on ,
219, 270, 380
influence of gravitational forces on ,
271, 344, 345
influence of magnetic forces on , 380
in processes involving radiation,
335345
role of internal energy in , 82, 376,
378
role of thermodynamic potential in ,
93
role of work function in , 92
Equilibrium conditions (explicit)
general statement of , 112, 113, 200,
379
in dilute solutions, 155157
in processes involving radiation,
335336
in terms of activities, 200, 201
Equilibrium constant, 139, 156, 200,
235, 236, 239, 336
Equipartition of energy, 79
Euler's theorem (about homogeneous
functions), 102, 104
Eutectic alloy, 188
Eutectic point, 188
ternary , 196
Exothermic process, 27
Expansion (thermal), 1
coefficient of , 3, 5
in vicinity of T = 0, 228, 258
Extensive quantities, 384
Extinction, atmospheric , 392, 393
402
SUBJECT INDEX
Fermi Dirac gas, see: degenerate perfect
gas
First law of thermodynamics, 2751
formulation of , 3739
history of , 2734.
Fluctuations, 389, 390
of density, 390392
Free energy, see: work function and
thermodynamic potential
Freezing point (see also: fusion), 116
^"ugacity, 197
of perfect gas, 197
of Van der Waals gas, 198
Fundamental point (see also:
triple point), 110, 126128
Fundamental state, 110, 126128
Fusion, point of , 116, 119
lowering of due to solutes, 166,
167, 202, 203
Fusion, process of , 219
Gadolinium sulphate, specific heat of ,
327
magnetization of , 349
Galvanic cells, 94, 95, 287, 288
electronomotive force of , 94, 289, 290
Galvanomagnetic effects, 368
Gas constant, 6, 7,398
Gas mixtures, 7, 13
activities of binary , 205
chemical equilibrium in , 137147,
236239, 270, 284, 336
heat capacity in , 315318
velocity of sound in , 318321
Generic probability definition, 78, 246
connection of with principle of
indetermination, 260
GibbsHelmholtz equation, 93, 96, 225
Grammolecule, 6
Gravitational forces, influence of on
equilibrium, 271, 344
H
Heat, 25
latent , see: latent heat
of reaction, 27, 67
molal (see also: specific heat), 26
specific , see: specific heat
Heat capacity (see also: specific heat).
25,26
at constant magnetic field, 351, 352
at constant magnetization, 351, 352
at constant pressure, 40, 46, 65, 66
at constant volume, 40, 65, 66
of electron pairs, 388
of reacting gas mixtures, 315318
Heat content, see: heat function
Heat function, 46, 85, 86
connection of with latent heat and
heat of reaction, 67
of electron pairs, 388
of perfect gases, 47
role of in JouleThomson proc
ess, 69
Helium, curve of fusion and sublimation
in , 128
equilibrium of and hydrogen, 342
inversion point of , 72
supposed second order transformation
in 129
Henry's law, 159162
Heterogeneous system, 2
Homogeneous system, 2
Hydrogen, annihilation of , 343
equilibrium of and deuterium, 340
equilibrium of and helium, 342
solubility of in molten metals, 162
specific heat of , 44, 303
Hydrogen bromide, dissociation of ,
147
Hydrogen iodide, dissociation of , 144,
145
I
Image force, 273, 274
Independent component, 107110
systems with one , 115133
Independent variables, 98, 99, 107
Integrability, 22, 23
Integrating multiplier, 22, 58, 74
Intensitaet, 383, 384
Intensive quantities, 102, 384
Internal energy, 38, 83
constant of , 330
of degenerate perfect gases, 255
of dilute and perfect solutions,
151, 152
of electron gas, 278
SUBJECT INDEX
403
Internal energy, Continued
of perfect gas, 43
of Van der Waals gas, 65
Internal pressure, 9
Inversion temperature in JouleThomson
effect, 71, 72
Iodine, dissociation of , 148
distribution of between water and
carbon tetrachloride, 158
Ion clouds, 266287
lonization in solar atmosphere, 283285
Irreversible process, 35, 36
Isobaric process, 46, 47
Isochoric process, 47
Isosmotic solutions, 170
Isothermal process, 47, 48
work in , 87
Isothermals, of Van der Waals gas, 10
of CO 2 , 17
Isotonic, see: isosmotic
JouleThomson effect, 7073
differential , 70
integral , 73
JouleThomson process, 42
theory of , 7073
K
KamerlinghOnnes' empirical reduced
equation of state, 16
Kelvin's formula for influence of capil
larity on boiling pressure, 214
Kelvin's scale of temperatures, see:
absolute temperature
Lagrangean multipliers, 98
Langevin's theory of magnetization,
346349
LangevinBrillouin, magnetic equation
of state of , 348
Latent heat, 26, 46, 67, 118, 121
temperature dependence of , 121
LeChatelierBraun, principle of , 375
restricted principle of , 380383
Legendre transformation, 85
Lenz's principle of electrodynamics, 375
Liquidus curve, 185
Liquidus surface, 195
Lowering, of boiling pressure, 162167
of point of fusion, 166, 167, 202, 203
M
Magnetic equation of state, 347
of LangevinBrillouin, 348
Magnetics, 346
perfect , 348, 350, 353
Magnetization, 346
influence of on equilibrium, 380
Magnetocaloric cooling, 352354
Magnetostriction, 350
Mass fraction, 109, 182, 192, 193, 196
Mass law, 138140, 176178, 270, 271,
336
Mixture, 149
of gases, see: gas mixtures
Mobile equilibrium, principle of , 156
Mol,6
fraction, 13, 102, 136
Molal heat (see also: specific heat), 26
Molal volume, 6
of dilute and perfect solutions, 152
Multiplier, integrating , 22
Lagrangean , 98
N
NagaokaHonda effect, 350
Nernst's distribution law, 157, 158
Nernst's heat postulate, 225250
generality of , 246250
independence of from second law,
244
Nernst's own formulation of , 226,
227, 230
Planck's formulation of , 230, 231
Neumann's law, 45, 327
Neutrons, 260, 339, 343
equilibrium of and protons, 339
Nitric oxide, specific heat of , 44, 310
Nitrogen, solubility of in water, 161
specific heat of , 307, 308
Nitrogen tetroxide, dissociation of ,
141, 142
dissociation of in chloroform solu
tion, 177
heat capacity of , 319
Notations, 395
comparative table of , 396
404
SUBJECT INDEX
Opalescence in critical state, 393
Osmotic pressure, 167178
equation for , 169
Van t'Hoff s equation for in dilute
solutions, 171
OstwaldFreundlich, formula for solu
bility of small particles of , 218
Oxygen, specific heat of , 44, 307, 308
Partial molal quantities, 84
Partial molal thermodynamic potential,
90,91
in dilute solutions, 155, 200
in gases, 137
in strong electrolytes, 293
Peltier coefficient, 362
Perfect gases, 48
degenerate , 252265, 277, 278
mixtures of , 7, 134147
Perfect magnetics, 348, 350, 353
Perfect solutions, 151, 179
Perpetual motion machine of the second
kind, 54
Phase, 2
Phase equilibrium, 100, 101, 104114
displacement of , 156, 376380
in binary systems, 179193
in dilute solutions, 154167
in pure substances, 115128
number of phases in , 110, 111
of higher order, 128133
temperature and pressure in , 104
Phase rule, 110, 111
Phosphonium chloride, curve of fusion
and sublimation of , 128
Photoelectric effect, 275, 280
Piezoelectricity, 360, 361
Point of fusion, 116128
influence of capillarity on , 219
lowering of by solutes, 166, 167,
202, 203
Point of inversion, see: inversion point
Point of transition, see: transition point
Polarization, electric , 359, 360
Positrons (or positive electrons), 337
equilibrium of and negative elec
trons, 337, 338
Principle, Boltzmann's , 78, 79, 240,
260, 296, 385
entropy , 6062, 82
Lenz's of electrodynamics, 375
of indetermination, 260
of LeChatelierBraun, 375, 380383
of mobile equilibrium, 156
of ThomsenBerthelot, 226
Probability, 7679, 246, 247, 260
generic definition of 78, 246
relation of generic definition to prin
ciple of indetermination, 260
specific definition of , 78
Protons, annihilation of , 343, 344
equilibrium of and neutrons, 339
Pure substances, phase equilibrium in ,
115128
Pyroelectricity, 361
Q
Quadruple point, 179
Quantitaet, 383, 384
R
Radiation, black , 331
energy of , 332, 333
entropy of , 332, 333
pressure of , 332
RamsayShields, formula for surface
tension of , 221
Raoult's law, 162, 163, 164
Rayleigh's formula for atmospheric light
scattering, 392
Reciprocity relations, 21, 64, 70, 89,
350, 360, 381
Reduced variables, 12, 16, 68, 69, 71, 73,
123
Regelation, 120
Reversible process, 35
Richardson's formula for thermionic cur
rent, 276
Rigid envelope, 100, 101
SackurTetrode formula (for absolute
entropy of gases), 237241, 257
Saturated solution, 174, 176
Scale of temperatures, 1
absolute ,8,58, 7476
centrigrade , 5
SUBJECT INDEX
405
Solutions),
Scattering of light, 391, 392
Second law of thermodynamics, 53, 54,
58, 59, 62, 82
Semiconductors, 368374
thermoelectric effect in , 276, 365
Semipermeable membrane, 101, 135, 168
Sieverts' law (solubility of gases in molt
en metals), 162
Simon and Glatzel, formula for pressure
of fusion of t 122
Simple system, 2
Solidus curve, 185
Solidus surface, 195
Solubility, mutual (complete or par
tial), 149, 150, 184, 186, 189, 190
of gases, see: Henry's law
of solids in liquids, 174, 218
Solute, 151
Solution, 149151
dilute (see also:
152, 177
perfect ,151, T'
Solvent, 151 .ula for electronic
Sommerf eld's >f metals, 278, 326
specific heanula for inner potential
Sommerf eld's f'
of met ; ' ula for thermoelectric
Sommerf ela 068
power, 2&\ also: heat capacity),
Specific heat (s\
26, 296329 sure, 40, 46, 65, 66
at constanprefc. 4Q ( 55 ? ^
atconstajvolumct^ 327
doublet , >1, 310, 3 ?2 p
in vicinr f T Q t
of elect <ytes, 329
of elect" 18 i n metals, 278, 326
of gas*, 44, 298311
of liqi*s, 328, 329
of pei** gases, 44
of sois, 45, 233, 321327
of suites, 329
rotaticA 1 ~ 29 ^ 300, 304, 305, 307
transl/onal ,299
vibraanal , 302, 303, 307309, 311
Specifitf> roDa bih'ty definition, 78
Specifi. quantities and properties, 102,
3/
Stabi# conditions of dynamical , 11,
4
Stability, Continued
conditions of thermodynamical ,
117,131,381
Statistical
average, 76
point of view, 7679, 246248,
385, 386
principles, 7679, 259, 260, 296
Strong electrolytes, 204, 287
activity function of ^ 204 205, 294
theory of , 29029* 361
Sublimation, 126, lX
Sugar solution, '*notic pressure of ,
172
s . states, 297, 300304
^Jpercooled liquids, 248
Superheated state, 11, 117, 126
Supraconductive transition points, 355
Supraconductivity, 354358
Surface tension, 209224
influence of on boiling and melting
points, 214, 216219
temperature dependence of , 224
Symmetry number, 300, 313, 314
System, thermodynamical 2
Table
comparative 1 of notations, 396
 of conv ersion factorSf 397
of f f andamental constants, 398
ratine, 1
a)F ,olute ,8,58, 7476
centrigrade , 5
Tension, coefficient of , 3
Ternary systems, 194196
Thermal equation of state (see: equation
of state), 3
Thermal equilibrium, see: equilib
rium
Thermal expansion (see also: expansion)
coefficient of , 3
in vicinity of T 0, 228
Thermal properties of matter, 27
Thermionic emission, 274
in classical theory, 279
in quantum theory, 279
Thermionic work function, 274, 276, 279,
280, 282
406
SUBJECT INDEX
Thermodymanic potential, 90
connection of with nonmechanical
work, 90
of degenerate perfect gas, 257
of electron gas, 278
of gas mixtures, 137, 207
of perfect gas, 91, 331
of Van der Waals' gas, 91
role of in equilibrium, 9193, 106,
112, 113,115
Thermodynamics 1
limitations of , 387393
subject of , 2
Thermoelectric effect, 3oi <^%
empirical ,364
in classical theory, 273, 367
in quantum theory, 282, 368
Thermoelectric e.m.f., 368
Thermoelectric power, 364, 366, 368
Thermomagnetic effects, 368
Third law of thermodynamics, see:
Nernst's postulate
ThomsenBerthelot, principle of , 226
Thomson coefficient (in thermoelec
tricity), 362, 365, 366
Transformation pressure and tempera
ture, 116, 121, J22
in binary systems. 183193
01A 910
influence of capillarity 7 on ~~ Z1 *' '
influence of solutes on " 162 " 167
Transition points of supravr onductors '
355 __
influence of magnetic field J '
356358
Trichloroacetic acid, distribution of
between water and ethyl ether, 158
Triple point, 125128
Trouton and Deprez, rule of , 124,
125
U
Unattainability of absolute zero, 243245
Units, conversion factors of , 397
Van der Waals'
equation of state, 9, 12
equation of state for binary gas
mixtures, 13, 206
formula for vapor pressure, 121
formula for surface tension, 221
Van t'Hoff's equation for osmotic
pressure, 171
Vapor pressure, see: boiling pressure
in binary systems, 180, 182
'iations, 82
_ 'tv of sound, 49
_ j n Jegenerate gases, 264265

Virtual chan ixtures ' 318 ~ 321
Virtual dispfi 82 t t ,
9799 "nents, method of ,
Vv
Water, activity and f i
^9g of liquid ,
Water vapor, dissociatit
Weak electrolytes, 20*i of , 143
Work, 19  287
element of 18,
Work function, '**'^ 210
of perfect^ 7 '
of stron^ f 8 ' 8 ? \
s electrolytes 293
  of Van der Waals ga&R9
role of in equilibriumAl, 92
thermionic , see: ther^onic work
function