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Coulomb interaction; this is the case for the electron“phonon interaction, which originates in the
Coulomb interaction between the electrons and nuclei constituting a piece of material such as a
metal.
48 2. Operators on the multi-particle state space


The total electron“photon Hamiltonian, for the case of non-relativistic electrons,
then becomes
H = Hph + Hel + Hel’ph (2.79)
where
1
(P ’ eA(x))2
Hel + Hel’ph = (2.80)
2m
and P is the total momentum operator for the electrons, Eq. (2.14).
Show that the electron“photon interaction can be written in the form
e2
Hel’ph = ’ dx jp (x) · A(x) + dx n(x) A2 (x) (2.81)
2m
where the current and density operators for the electrons are speci¬ed in Sections 2.3
and 2.2.


2.5 The statistical operator
Up until now, we have described the physical states of a system in terms of state
vectors in the multi-particle state space. A general state vector, |Ψ , can be expanded
on the basis vectors (using for once the resolution of the identity on the multi-particle
state space)

1
|Ψ = p1 ∨ p2 ∨ · · · ∨ pN |Ψ |p1 ∨ p2 ∨ · · · ∨ pN (2.82)
N! p
1 ,...,pN
N =0

or expressed in terms of the vacuum state with the help of our new, so far only
kinematic gadget, the ¬eld operator

c(p1 , . . . , pN ) a† 1 · · · a† N |0 ,
|Ψ = (2.83)
p p
N =0 p1 ,...,pN

where the cs are the complex amplitude coe¬cients specifying the state. Equivalently
the state vector could be expressed in terms of the ¬eld operator in the position
representation.
The description of a system in terms of its wave function is not the generic one
as systems often can not be considered isolated, and instead will be in a mixture of
states described by a statistical operator or density matrix (at a certain moment in
time)
ρ≡ pk |ψk ψk | (2.84)
k
allowing only the statements for the events of the system that it occurs with prob-
ability pk in the quantum states |ψk in the multi-particle state.11 Certainly this is
11 In general a statistical operator is speci¬ed by a set of (normalized) state vectors, |ψ1 ,
|ψ2 , . . . , |ψn , not necessarily orthogonal, and a set of non-negative numbers adding up to one,
n n
i=1 pi |ψi ψi |. Since the statistical operator is hermitian and
i=1 pi = 1, according to ρ =
non-negative, it is always possible to ¬nd an orthonormal set of states |φ1 , |φ2 , . . . , |φN , so that
ρ = N πn |φn φn |, where πn ≥ 0 and N πn = 1.
n=1 n=1
2.5. The statistical operator 49


the generic situation in condensed matter physics and statistical physics in general.
A diagonal element of the statistical operator, ψ|ρ|ψ , thus gives the probability for
the occurrence of the arbitrary state |ψ .
For the evaluation of the average value of a physical quantity represented by the
operator A, a mixture adds an additional purely statistical averaging, as the quantum
average value in a state is weighted by the probability of occurrence of the state

A≡ pk ψk |A|ψk . (2.85)
k

In view of Eq. (2.84), the average value for a mixture can be expressed in terms of
the statistical operator according to

A = Tr(ρ A) (2.86)

where Tr denotes the trace in the multi-particle state, i.e. the sum of all diagonal
elements evaluated in an arbitrary basis, generalizing the matrix element formula for
the average value in a pure state |ψ , A = Tr(|ψ ψ| A).
The statistical operator is seen to be hermitian and positive, ψ| ρ |ψ ≥ 0, and
has unit trace for a normalized probability distribution. The statistical operator is
only idempotent, i.e. a projector, for the case of a pure state, ρ = |ψ ψ|. For a
mixture we have ρ2 = ρ, and Tr ρ2 < 1.
In practice the most important mixture of states will be the one corresponding to
a system in thermal equilibrium at a temperature T (including as a limiting case the
zero temperature situation where the system de¬nitely is in its ground state). In that
case, applying the zeroth law of thermodynamics (that two systems in equilibrium
at temperature T will upon being brought in thermal contact be in equilibrium at
the same temperature) gives, for the thermal equilibrium statistical operator (Boltz-
mann™s constant, the converter between energy and absolute temperature scale, is
denoted k)
1 ’H/kT
ρT = e , (2.87)
Z
where the normalization factor

Z(T, V, Ns ) = Tr e’H/kT = e’F (T,V,Ns )/kT (2.88)

expressing the normalization of the probability distribution of Boltzmann weights,
is the partition function. Here Tr denotes the trace in the multi-particle state space
of the physical system of interest, but since the number of particles is ¬xed there is
only the contributions from the corresponding N -particle subspace

e’H/kT = e’En /kT (2.89)
n

as we are describing the system in the canonical ensemble. As the temperature ap-
proaches absolute zero, the term with minimum energy, the non-degenerate ground
state energy, dominates the sum, and at zero temperature, the average value of a
physical quantity becomes the N -particle ground state average value. The partition
50 2. Operators on the multi-particle state space


function or equivalently the free energy, F , are functions of the macroscopic param-
eters, the temperature, T , and the volume, V , and the number of particles, Ns , of
di¬erent species in the system, and it contains all thermodynamic information.
When a system consists of a huge number of particles, it is more convenient to
perform calculations in the grand canonical ensemble, where instead of the inconve-
nient constrain of a ¬xed number of particles, their chemical potential and average
number of particles are speci¬ed. In the grand canonical ensemble, the system is thus
described in the multi-particle state space. The system is thus imagined coupled to
particle reservoirs, or a subsystem is considered. The system can exchange particles
with the reservoirs which are described by their chemical potentials (assuming in
general several particle species present). This feature is simply included by introduc-
ing Lagrange multipliers, i.e. tacitly understanding that single-particle energies are
measured relative to their chemical potential, H ’ H ’ s μs Ns . In this case, the
partition function instead of being a function of the number of particles, is a function
of the chemical potentials for the species in question

Zgr (T, μs ) = Tr e’(H’μs Ns )/kT = e’©(T,μs )/kT (2.90)

speci¬ed by the average number of particles according to

‚©
Ns = ’ . (2.91)
‚μs T,V

For systems of particles where the total number of particles is not conserved,
i.e. where the Hamiltonian and the total number operator do not commute, such as
for phonons and photons, the chemical potential vanishes, and the grand canonical
ensemble is employed. For degenerate fermions, such as electrons in a metal, the
chemical potential is a huge energy compared to relevant temperatures, viz. tied to
the Fermi energy as discussed in Exercise 2.15.
The average value in the grand canonical ensemble of a physical quantity repre-
sented by the operator A is thus

e(©’En +μN )/kT N, En |A|En , N .
A = Tr(ρ A) = (2.92)
N,n

As the following exercise shows, if alternatively attempted in the canonical ensem-
ble, calculations run smoothly in the grand canonical ensemble free of the constraint
of a ¬xed number of particles. In the thermodynamic limit, using either of course
gives the same results as the ¬‚uctuations in the particle number in the grand canon-
ical ensemble around the mean value then is negligible.

Exercise 2.14. Show that the grand canonical partition function for non-interacting
non-relativistic fermions or bosons of mass m contained in a volume V is given by

p2
ln (1 “ e’( p ’μ)/kT
) = e’©0 /kT
ln Zgr (T, V, μ) = “
(0)
, = (2.93)
p
2m
p
2.5. The statistical operator 51


and the average number of particles are speci¬ed by the Bose“Einstein or Fermi“Dirac
distributions, respectively,
‚©0 1
=’
N = (2.94)
p ’μ)/kT “1
e(
‚μ p
T,V

from which one readily veri¬es that the thermodynamic potential is speci¬ed by the
pressure, P , and volume of the system according to
©0 (T, μ, V ) = ’P V . (2.95)
Exercise 2.15. Show that for a system of non-interacting degenerate fermions, i.e.
at temperatures where kT μ, the chemical potential is pinned to the Fermi energy,
2 2 2/3
F= (3π n) /2m, as
2
T
1’a
μ= , (2.96)
F
TF
where the fermions of mass m are assumed residing in three spatial dimensions (in
which case the constant of order one is a = π 2 /12) with a density n, and TF = F /k
is the Fermi temperature which for a metal, in view of the large density of conduction
electrons, is seen to be huge, typically of the order 104 K.
Exercise 2.16. At zero temperature, a system of fermions such as a metal contains
high-speed electrons, all states below the Fermi energy are fully occupied, a reservoir
for injecting electrons into other conductors. For bosons the opposite, coming to rest,
can happen. First we observe that for non-interacting bosons the chemical potential
can not be positive, μ ¤ 0, as dictated by the Bose“Einstein distribution function
for occupation of energy levels. As the temperature decreases, the chemical potential
increases and becomes vanishingly small at and below the temperature T0 determined
by the density, n, and mass, m, of the bosons (say, spinless and enclosed in a volume
V ) according to

(m)3/2 1/2
N
=√
n= d . (2.97)
’1
/kT0
V e
2 π2 3
0
At this temperature, the lowest energy level, p=0 = 0, starts to be macroscopically
occupied.
Show that at temperatures below T0 , the number of bosons in the lowest level is
(the population of the other levels are governed by the Bose“Einstein distribution)
3/2
T
1’
N0 = N . (2.98)
T0
In the degenerate region at temperatures below T0 , the bosons comes to rest, the
phenomenon of Bose“Einstein condensation (1925), the bosons become ordered in
momentum space. The total condensation at zero temperature is of course a trivial
feature of the quantum statistics of non-interacting bosons. Using the model of non-
interacting bosons to estimate T0 for the case of 4 He gives T0 3.2 K, quite close to
the temperature 2.2 K of the »-transition where liquid Helium becomes a super¬‚uid
(discovered 1928 and proposed a Bose“Einstein condensate by Fritz London, 1938).
52 2. Operators on the multi-particle state space


2.6 Summary
In this chapter we have constructed the operators of relevance on the multi-particle
space, and shown how they are expressed in terms of the quantum ¬elds, the creation
and annihilation ¬elds. The kinematics of a many-body system, its possible quantum
states and the operators representing its physical quantities, is thus expressed in
terms of these two objects. The Hamiltonians on the multi-particle state space were
constructed for the generic types of many-body interactions. We now turn to consider
the dynamics of many-body systems described by their quantum ¬elds on a multi-
particle state space. In particular the quantum dynamics of a quantum ¬eld theory
describing systems out of equilibrium. This will lead us to the study of correlation
functions for quantum ¬elds, the Green™s functions for non-equilibrium states.
3

Quantum dynamics and
Green™s functions

In the previous chapter we studied the kinematics of many-body systems, and the
form of operators representing the physical properties of a system, all of which were
embodied by the quantum ¬eld. In this chapter we shall study the quantum dynamics
of many-body systems, which can also be embodied by the quantum ¬elds. We shall
employ the fact that the quantum dynamics of a system, instead of being described
in terms of the dynamics of the states or the evolution operator, i.e. as previously
done through the Schr¨dinger equation, can instead be carried by the quantum ¬elds.
o
The quantum dynamics is therefore expressed in terms of the correlation functions
or Green™s functions of the quantum ¬elds evaluated with respect to some state of
the system. In particular we shall consider the general case of quantum dynamics
for arbitrary non-equilibrium states. After introducing various types of Green™s func-
tions and relating them to measurable quantities, we will discuss the simpli¬cations
reigning for the special case of equilibrium states.


3.1 Quantum dynamics
Quantum dynamics can be described in di¬erent ways since quantum mechanics is
a linear theory and the dynamics described by a unitary transformation of states.1
This will come in handy in the next chapter when we study a quantum theory in
terms of its perturbative expansion using the so-called interaction picture. Here we
¬rst discuss the Schr¨dinger and Heisenberg pictures.
o
1 This should be contrasted with classical mechanics where dynamics is speci¬ed in terms of
the physical quantities themselves, the generic case being intractable nonlinear partial di¬erential
equations. We shall study such a classical situation in Chapter 12 with the help of methods borrowed
from quantum ¬eld theory, and where in addition the classical system interacts with an environment
as described by a stochastic force.




53
54 3. Quantum dynamics and Green™s functions


3.1.1 The Schr¨dinger picture
o
Having the Hamiltonian on the multi-particle space at hand we can consider the
dynamics described in the multi-particle space. An arbitrary state in the multi-
particle space has at any time in question the expansion on, say the position basis
states

|Ψ(t) dx1 dx2 . . . dxN ψN (x1 , x2 , . . . , xN , t) |x1 3x2 · · · 3xN ,
= (3.1)
N =0


where 3 stands for ∨ or § for the bose or fermi cases respectively. The coe¬cients

ψN (x1 , x2 , . . . , xN , t) = x1 3x2 · · · 3xN |Ψ(t) (3.2)

are the wave functions describing each N -particle system, and they are symmetric
or antisymmetric due to the symmetry properties of the basis states |x1 3x2 · · · 3xN ,
i.e. no new state is produced by using non-symmetric coe¬cients ψN (x1 , x2 , . . . , xN , t).
The dynamics of a multi-particle particle state is speci¬ed by the Schr¨dinger
o
equation in the multi-particle space
d|Ψ(t)
= H(t) |Ψ(t) ,
i (3.3)
dt
where H(t) is the Hamiltonian on the multi-particle space, which can be explic-
itly time dependent due to external forces as our interest will be to consider non-
equilibrium states. In the multi-particle space, the dynamics of all N -particle sys-
tems are described simultaneously since the above equation contains the in¬nite set
of equations, N = 0, 1, 2, . . ., which in the position representation are

‚ψN (x1 , x2 , . . . , xN , t)
i = dx1 dx2 . . . dxM ψM (x1 , x2 , . . . , xM , t)
‚t
M=0


— x1 3x2 · · · 3xN |H(t)|x1 3x2 · · · 3xM . (3.4)

The even or odd character of a wave function is preserved in time as any Hamil-
ˆ ˆ
tonian for identical particles is symmetric in the pi s and xi s as well as other degrees
of freedom (this is the meaning of identity of particles, no interaction can distinguish
them), so if even- or oddness of a wave function is the state of a¬airs at one moment
in time it will stay this way for all times.2
For the case of two-particle interaction, Eq. (2.59), the Hamiltonian has only
nonzero matrix elements between con¬gurations with the same number of particles,
the total number of particles is a conserved quantity, and the in¬nite set of equa-
tions, Eq. (3.4), splits into independent equations describing systems with the de¬nite
number of particles N = 0, 1, 2, . . . .3 For N = 0, the vacuum state, we have for the
2 Time-invariant subspaces other than the symmetric and anti-symmetric ones do not seem to be
physically relevant.
3 For the case of an N -particle system, the multi-particle space is then not needed, we could

discuss it solely in terms of its N -particle state space.
3.1. Quantum dynamics 55


c-number representing its wave function
dψ0 (t)
i = 0|H(t)|0 ψ0 (t) = 0 , (3.5)
dt
where the last equality sign follows from the fact that since the Hamiltonian for two-
particle interaction, Eq. (2.59), operates ¬rst with the annihilation operator on the
vacuum state, it annihilates it. Since Hamiltonians are normal-ordered, this result is
quite general: the vacuum state is without dynamics.
In the case of electron“phonon interaction, the Hamiltonian has o¬-diagonal ele-
ments with respect to the phonon multi-particle subspace. The number of phonons is
owing to interaction with the electrons not conserved; an electron can emit or absorb
phonons just like an electron in an excited state of an atom can emit a photon in the
decay to a lower energy state. The chemical potential of phonons thus vanishes.
Instead of describing the dynamics in terms of the state vector we can introduce
the time development or time evolution operator4
|ψ(t) = U (t, t ) |ψ(t ) (3.6)

connecting the state vectors at the di¬erent times in question where they provide a
complete description of the system. Solving the Schr¨dinger equation, Eq. (3.3), by
o
iteration gives
U (t, t ) = T e’ t dt H(t) ,
t¯ ¯
i
(3.7)
where T denotes time-ordering.5 The time-ordering operation orders a product of
time-dependent operators into its time-descending sequence (displayed here for the
case of three operators)
§
⎨ A(t1 ) B(t2 ) C(t3 ) for t1 > t2 > t3
B(t2 ) A(t1 ) C(t3 ) for t2 > t1 > t3
T (A(t1 ) B(t2 ) C(t3 )) = (3.8)
©
etc. etc.

In case of fermi ¬elds, the time-ordering (and anti-time-ordering which we shortly en-
counter) shall by de¬nition include a product of minus signs, one for each interchange
of fermi ¬elds. Since the Hamiltonian contains an even number of fermi ¬elds, no sta-
tistical factors are thus involved in interchanging Hamiltonians referring to di¬erent
moments in time under the time-ordering symbol.
The Schr¨dinger equation, Eq. (3.3), then gives the equation of motion for the
o
time evolution operator6
‚U (t, t )
= H(t) U (t, t ) .
i (3.9)
‚t
4 It is amazing how compactly quantum dynamics can be captured, encapsulated in the single
object, the evolution operator.
5 For details see, for example, chapter 2 of reference [1].
6 From a mathematical point of view, convergence properties of limiting processes for operator

sequences At are inherited from the topology of the vector space; i.e. convergence is de¬ned by
convergence of an arbitrary vector At |ψ . The dual space of linear operator on the multi-particle
state space can also be equipped with its own topology by introducing the scalar product for
operators A and B, Tr(B † A). But the result of di¬erentiating is gleaned immediately from simple
algebraic properties.
56 3. Quantum dynamics and Green™s functions

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