Chapter 4

Introduction to many-body
quantum mechanics

4.1 The complexity of the quantum many-body prob-

lem

After learning how to solve the 1-body Schrödinger equation, let us next generalize to
more particles. If a single body quantum problem is described by a Hilbert space H
of dimension dimH = d then N distinguishable quantum particles are described by the
tensor product of N Hilbert spaces

N
O
H(N ) ≡ H⊗N ≡ H (4.1)
i=1

with dimension dN .
As a first example, a single spin-1/2 has a Hilbert space H = C2 of dimension 2,
N
but N spin-1/2 have a Hilbert space H(N ) = C2 of dimension 2N . Similarly, a single
particle in three dimensional space is described by a complex-valued wave function ψ(~x)
of the position ~x of the particle, while N distinguishable particles are described by a
complex-valued wave function ψ(~x1 , . . . , ~xN ) of the positions ~x1 , . . . , ~xN of the particles.
Approximating the Hilbert space H of the single particle by a finite basis set with d
basis functions, the N-particle basis approximated by the same finite basis set for single
particles needs dN basis functions.
This exponential scaling of the Hilbert space dimension with the number of particles
is a big challenge. Even in the simplest case – a spin-1/2 with d = 2, the basis for N = 30
spins is already of of size 230 ≈ 109 . A single complex vector needs 16 GByte of memory
and will not fit into the memory of your personal computer anymore.

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 This challenge will be to addressed later in this course by learning about

1. approximative methods, reducing the many-particle problem to a single-particle

problem

2. quantum Monte Carlo methods for bosonic and magnetic systems

3. brute-force methods solving the exact problem in a huge Hilbert space for modest
numbers of particles

4.2 Indistinguishable particles

4.2.1 Bosons and fermions
In quantum mechanics we assume that elementary particles, such as the electron or
photon, are indistinguishable: there is no serial number painted on the electrons that
would allow us to distinguish two electrons. Hence, if we exchange two particles the
system is still the same as before. For a two-body wave function ψ(~q1 , ~q2 ) this means
that
ψ(~q2 , ~q1 ) = eiφ ψ(~q1 , ~q2 ), (4.2)
since upon exchanging the two particles the wave function needs to be identical, up to
a phase factor eiφ . In three dimensions the first homotopy group is trivial and after
doing two exchanges we need to be back at the original wave function1

ψ(~q1 , ~q2 ) = eiφ ψ(~q2 , ~q1 ) = e2iφ ψ(~q1 , ~q2 ), (4.3)

and hence e2iφ = ±1:

ψ(~q2 , ~q1 ) = ±ψ(~q1 , ~q2 ) (4.4)
The many-body Hilbert space can thus be split into orthogonal subspaces, one in which
particles pick up a − sign and are called fermions, and the other where particles pick
up a + sign and are called bosons.

Bosons
For bosons the general many-body wave function thus needs to be symmetric under
permutations. Instead of an arbitrary wave function ψ(~q1 , . . . , ~qN ) of N particles we
use the symmetrized wave function
X
Ψ(S) = S+ ψ(~q1 , . . . , ~qN ) ≡ NS ψ(~qp(1) , . . . , ~qp(N ) ), (4.5)
p

where the sum goes over all permutations p of N particles, and NS is a normalization
factor.
1
As a side remark we want to mention that in two dimensions the first homotopy group is Z and not
trivial: it matters whether we move the particles clock-wise or anti-clock wise when exchanging them,
and two clock-wise exchanges are not the identity anymore. Then more general, anyonic, statistics are
possible.

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Fermions
For fermions the wave function has to be antisymmetric under exchange of any two
fermions, and we use the anti-symmetrized wave function
X
Ψ(A) S− ψ(~q1 , . . . , ~qN ) ≡ NA sgn(p)ψ(~qp(1) , . . . , ~qp(N ) ), (4.6)
p

where sgn(p) = ±1 is the sign of the permutation and NA again a normalization factor.
A consequence of the antisymmetrization is that no two fermions can be in the same
state as a wave function
ψ(~q1 , ~q2 ) = φ(~q1 )φ(~q2 ) (4.7)
since this vanishes under antisymmetrization:

Ψ(~q1 , ~q2 ) = ψ(~q1 , ~q2 ) − ψ(~q2 , ~q1 ) = φ(~q1 )φ(~q2 ) − φ(~q2 )φ(~q1 ) = 0 (4.8)

Spinful fermions
Fermions, such as electrons, usually have a spin-1/2 degree of freedom in addition
to their orbital wave function. The full wave function as a function of a generalized
coordinate ~x = (~q, σ) including both position ~q and spin σ.

4.2.2 The Fock space

The Hilbert space describing a quantum many-body system with N = 0, 1, . . . , ∞
particles is called the Fock space. It is the direct sum of the appropriately symmetrized
single-particle Hilbert spaces H:
M∞
S± H⊗n (4.9)
N =0

where S+ is the symmetrization operator used for bosons and S− is the anti-symmetrization
operator used for fermions.

The occupation number basis

Given a basis {|φ1 i, . . . , |φL i} of the single-particle Hilbert space H, a basis for the
Fock space is constructed by specifying the number of particles ni occupying the single-
particle wave function |f1 i. The wave function of the state |n1 , . . . , nL i is given by the
appropriately symmetrized and normalized product of the single particle wave functions.
For example, the basis state |1, 1i has wave function

1
√ [φ1 (~x1 )φ2 (~x2 ) ± φ1 (~x2 )φ2 (~x1 )] (4.10)
2
where the + sign is for bosons and the − sign for fermions.
For bosons the occupation numbers ni can go from 0 to ∞, but for fermions they
are restricted to ni = 0 or 1 since no two fermions can occupy the same state.

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The Slater determinant
The antisymmetrized and normalized product of N single-particle wave functions φi
can be written as a determinant, called the Slater determinant

N φ1 (~x1 ) · · · φN (~x1 )
Y 1 .. ..
S− φi (~xi ) = √ . . . (4.11)
N
i1 φ1 (~xN ) · · · φN (~xN )

Note that while the set of Slater determinants of single particle basis functions forms
a basis of the fermionic Fock space, the general fermionic many body wave function is a
linear superposition of many Slater determinants and cannot be written as a single Slater
determinant. The Hartee Fock method, discussed below, will simplify the quantum
many body problem to a one body problem by making the approximation that the
ground state wave function can be described by a single Slater determinant.

4.2.3 Creation and annihilation operators

Since it is very cumbersome to work with appropriately symmetrized many body wave
functions, we will mainly use the formalism of second quantization and work with
creation and annihilation operators.
The annihilation operator ai,σ associated with a basis function |φi i is defined as the
result of the inner product of a many body wave function |Ψi with this basis function
|φi i. Given an N-particle wave function |Ψ(N ) i the result of applying the annihilation
operator is an N − 1-particle wave function |Ψ̃(N ) i = ai |Ψ(N ) i. It is given by the
appropriately symmetrized inner product
Z
Ψ̃(~x1 , . . . , ~xN −1 ) = S± d~xN fi† (~xN )Ψ(~x1 , . . . , ~xN ). (4.12)

Applied to a single-particle basis state |φj i the result is

ai |φj i = δij |0i (4.13)

where |0i is the “vacuum” state with no particles.

The creation operator a†i is defined as the adjoint of the annihilation operator ai .
Applying it to the vacuum “creates” a particle with wave function φi :

|φi i = a†i |0i (4.14)

For sake of simplicity and concreteness we will now assume that the L basis functions
|φi i of the single particle Hilbert space factor into L/(2S + 1) orbital wave functions
fi (~q) and 2S + 1 spin wave functions |σi, where σ = −S, −S + 1, ..., S. We will write
creation and annihilation operators a†i,σ and ai,σ where i is the orbital index and σ the
spin index. The most common cases will be spinless bosons with S = 0, where the spin
index can be dropped and spin-1/2 fermions, where the spin can be up (+1/2) or down
(−1/2).

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Commutation relations
The creation and annihilation operators fulfill certain canonical commutation relations,
which we will first discuss for an orthogonal set of basis functions. We will later gener-
alize them to non-orthogonal basis sets.
For bosons, the commutation relations are the same as that of the ladder operators
discussed for the harmonic oscillator (2.62):

[ai , aj ] = [a†i , a†j ] = 0 (4.15)

[ai , a†j ] = δij . (4.16)

For fermions, on the other hand, the operators anticommute

{a†jσ′ , aiσ } = {a†iσ , ajσ′ } = δσσ′ δij

{aiσ , ajσ′ } = {a†iσ , a†jσ′ } = 0. (4.17)

The anti-commutation implies that

(a†i )2 = a†i a†i = −a†i a†i (4.18)

and that thus

(a†i )2 = 0, (4.19)
as expected since no two fermions can exist in the same state.

Fock basis in second quantization and normal ordering

The basis state |n1 , . . . , nL i in the occupation number basis can easily be expressed in
terms of creation operators:
L
Y
|n1 , . . . , nL i = (a†i )ni |0i = (a†1 )n1 (a†2 )n2 · · · (a†L )nL |0i (4.20)
i=1

For bosons the ordering of the creation operators does not matter, since the operators
commute. For fermions, however, the ordering matters since the fermionic creation
operators anticommute: and a†1 a†2 |0i = −a†1 a†2 |0i. We thus need to agree on a specific
ordering of the creation operators to define what we mean by the state |n1 , . . . , nL i.
The choice of ordering does not matter but we have to stay consistent and use e.g. the
convention in equation (4.20).
Once the normal ordering is defined, we can derive the expressions for the matrix
elements of the creation and annihilation operators in that basis. Using above normal
ordering the matrix elements are
Pi−1
ni
ai |n1 , . . . , ni , . . . , nL i = δni ,1 (−1) j=1 |n1 , . . . , ni − 1, . . . , nL i (4.21)
Pi−1
a†i |n1 , . . . , ni , . . . , nL i = δni ,0 (−1) j=1 ni |n1 , . . . , ni + 1, . . . , nL i (4.22)

where the minus signs come from commuting the annihilation and creation operator to
the correct position in the normal ordered product.

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4.2.4 Nonorthogonal basis sets
In simulating the electronic properties of atoms and molecules below we will see that
the natural choice of single particle basis functions centered around atoms will nec-
essarily give a non-orthogonal set of basis functions. This is no problem, as long as
the definition of the annihilation and creation operators is carefully generalized. For
this generalization it will be useful to introduce the fermion field operators ψσ† (~r) and
ψσ (~r), creating and annihilating a fermion localized at a single point ~r in space. Their
commutation relations are simply

{ψσ† ′ (~r), ψσ (r~′ )} = {ψσ† (~r), ψσ′ (r~′ )} = δσσ′ δ(~r − r~′ )
{ψσ (~r), ψσ′ (r~′ )} = {ψσ† (~r), ψσ† ′ (r~′ )} = 0. (4.23)

The scalar products of the basis functions define a matrix

Z
Sij = d3~rfi∗ (~r)fj (~r), (4.24)

which is in general not the identity matrix. The associated annihilation operators aiσ
are again defined as scalar products
X Z
aiσ = (S )ij d3~rfj∗ (~r)ψσ (~r).
−1
(4.25)
j

The non-orthogonality causes the commutation relations of these operators to differ

from those of normal fermion creation- and annihilation operators:

{a†iσ , ajσ′ } = δσσ′ (S −1 )ij

{aiσ , ajσ′ } = {a†iσ , a†jσ′ } = 0. (4.26)

Due to the non-orthogonality the adjoint a†iσ does not create a state with wave function
fi . This is done by the operator â†iσ , defined through:
X
â†iσ = Sji a†iσ , (4.27)
j

which has the following simple commutation relation with ajσ :

{â†iσ , ajσ } = δij . (4.28)

The commutation relations of the â†iσ and the âjσ′ are:

{â†iσ âjσ′ } = δσσ′ Sij

{âiσ , âjσ′ } = {â†iσ , â†jσ′ } = 0. (4.29)

We will need to keep the distinction between a and â in mind when dealing with
non-orthogonal basis sets.

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