8.9. Density Functional Theory (DFT)

8.9.1. Many Body Schrödinger Equation

We use (Hartree) atomic units in this whole section about DFT. We use the Born-Oppenheimer approximation, which says that the nuclei of the treated atoms are seen as fixed. A stationary electronic state (for N electrons) is then described by a wave function \Psi({\bf r_1},{\bf r_2},\cdots,{\bf r}_N) fulfilling the many-body Schrödinger equation

\hat H\ket\Psi=(\hat T+\hat U+\hat V)\ket\Psi=E\ket\Psi

where

\hat T = \sum_i^N -{1\over2}\nabla_i^2

is the kinetic term,

\hat U = \sum_{i<j}U({\bf r_i},{\bf r_j})= {1\over2}\sum_{i,j}U({\bf r_i},{\bf r_j})

U({\bf r_i},{\bf r_j})=U({\bf r_j},{\bf r_i})={1\over|{\bf r_i}-{\bf r_j}|}

is the electron-electron interaction term and

\hat V = \sum_i^N v({\bf r_i})

v({\bf r_i})=\sum_k -{Z_k\over|{\bf r_i}-{\bf R_k}|}

is the interaction term between electrons and nuclei, where R_k are positions of nuclei and Z_k the number of nucleons in each nucleus (we are using atomic units). So for one atomic calculation with the atom nucleus in the origin, we have just v({\bf r_i})=-{Z\over|{\bf r_i}|}.

|\Psi|^2=\Psi^*\Psi gives the probability density of measuring the first electron at the position \bf r_1, the second at \bf r_2, dots and the Nth electron at the position {\bf r}_N. The normalization is such that \int|\Phi|^2\d^3 r_1\d^3 r_2\dots\d^3 r_N=1. The \Psi is antisymmetric, i.e. \Psi({\bf r_1},{\bf r_2},\cdots,{\bf r}_N)= -\Psi({\bf r_2},{\bf
r_1},\cdots,{\bf r}_N)= -\Psi({\bf r_1},{\bf r}_N,\cdots,{\bf r_2}) etc.

Integrating |\Psi|^2 over the first N-1 electrons is the probability density that the N-th electron is at the position {\bf r}_N. Thus the probability density n({\bf r}) that any of the N electrons (i.e the first, or the second, or the third, dots, or the N-th) is at the position \bf r is called the particle (or number) density and is therefore given by:

n({\bf r})= \int \Psi^*({\bf r},{\bf r}_2,\cdots,{\bf r}_N) \Psi ({\bf r},{\bf r}_2,\cdots,{\bf r}_N) \,\d^3 r_2\,\d^3 r_3\cdots\d^3 r_N+

+\int \Psi^*({\bf r}_1,{\bf r},\cdots,{\bf r}_N) \Psi ({\bf r}_1,{\bf r},\cdots,{\bf r}_N) \,\d^3 r_1\,\d^3 r_3\cdots\d^3 r_N+\cdots

+\int \Psi^*({\bf r}_1,{\bf r}_2,\cdots,{\bf r}) \Psi ({\bf r}_1,{\bf r}_2,\cdots,{\bf r}) \,\d^3 r_1\,\d^3 r_2\,\d^3 r_3\cdots\d^3 r_{N-1}=

=\int(\delta({\bf r}-{\bf r}_1)+\delta({\bf r}-{\bf r}_2)+\cdots+\delta({\bf r}-{\bf r}_N))

\Psi^*({\bf r}_1,{\bf r}_2,\cdots,{\bf r}_N) \Psi ({\bf r}_1,{\bf r}_2,\cdots,{\bf r}_N) \,\d^3 r_1\,\d^3 r_2\,\d^3 r_3\cdots\d^3 r_{N}=

=\sum_{i=1}^N\int \braket{\Psi|{\bf r}_1,{\bf r}_2,\cdots,{\bf r}_N}\delta({\bf r}-{\bf r}_i) \braket{{\bf r}_1,{\bf r}_2,\cdots,{\bf r}_N|\Psi} \,\d^3 r_1\,\d^3 r_2\,\d^3 r_3\cdots\d^3 r_{N}=

=N\int \braket{\Psi|{\bf r}_1,{\bf r}_2,\cdots,{\bf r}_N}\delta({\bf r}-{\bf r}_1) \braket{{\bf r}_1,{\bf r}_2,\cdots,{\bf r}_N|\Psi} \,\d^3 r_1\,\d^3 r_2\,\d^3 r_3\cdots\d^3 r_{N}=

(8.9.1.1)=N\int \Psi^*({\bf r},{\bf r}_2,\cdots,{\bf r}_N) \Psi ({\bf r},{\bf r}_2,\cdots,{\bf r}_N) \,\d^3 r_2\,\d^3 r_3\cdots\d^3 r_N

Thus \int_\Omega n({\bf r})\,\d^3r gives the number of particles in the region of integration \Omega. Obviously \int
n({\bf r})\,\d^3r=N.

Note that the number density n({\bf r}) and potential V({\bf r}) in the Schroedinger equation is related to the electron charge density \rho({\bf r}) and electrostatic potential energy \phi({\bf r}) by:

\rho({\bf r}) = q n({\bf r})

q\phi({\bf r}) = V({\bf r})

where q is the particle elementary charge, which for electrons is q=-e=-1 in atomic units. The amount of electronic charge in the region \Omega is given by:

Q
    = \int_\Omega \rho({\bf r})\,\d^3r
    = q\int_\Omega n({\bf r})\,\d^3r
    = -\int_\Omega n({\bf r})\,\d^3r

The energy of the system is given by

(8.9.1.2)E=\braket{\Psi|\hat H|\Psi}= \braket{\Psi|\hat T|\Psi}+\braket{\Psi|\hat U|\Psi}+\braket{\Psi|\hat V|\Psi}= T+U+V

where

T=\braket{\Psi|\hat T|\Psi}=\sum_i^N\int \Psi^*({\bf r_1},{\bf r_2},\cdots,{\bf r_N})(-\half\nabla_i^2) \Psi({\bf r_1},{\bf r_2},\cdots,{\bf r_N})\,\d^3 r_1\,\d^3 r_2\cdots\d^3 r_N

U=\braket{\Psi|\hat U|\Psi}

V=\braket{\Psi|\hat V|\Psi}=\sum_i^N\int \Psi^*({\bf r_1},{\bf r_2},\cdots,{\bf r_N})v({\bf r_i}) \Psi({\bf r_1},{\bf r_2},\cdots,{\bf r_N})\,\d^3 r_1\,\d^3 r_2\cdots\d^3 r_N=

=\sum_i^N\int \Psi^*({\bf r_1},{\bf r_2},\cdots,{\bf r_N})v({\bf r_1}) \Psi({\bf r_1},{\bf r_2},\cdots,{\bf r_N})\,\d^3 r_1\,\d^3 r_2\cdots\d^3 r_N=

=N\int \Psi^*({\bf r_1},{\bf r_2},\cdots,{\bf r_N})v({\bf r_1}) \Psi({\bf r_1},{\bf r_2},\cdots,{\bf r_N})\,\d^3 r_1\,\d^3 r_2\cdots\d^3 r_N=

(8.9.1.3)=\int v({\bf r}) n({\bf r})\d^3 r=V[n]

It needs to be stressed, that E generally is not a functional of n alone, only the V[n] is. In the next section we show however, that if the \ket{\Psi} is a ground state (of any system), then E becomes a functional of n.

8.9.2. The Hohenberg-Kohn Theorem

The Schrödinger equation gives the map

C: V \to \Psi

where \Psi is the ground state. C is bijective (one-to-one correspondence), because to every V we can compute the corresponding \Psi from Schrödinger equation and two different V and V' (differing by more than a constant) give two different \Psi, because if V and V' gave the same \Psi, then by substracting

\hat H\ket{\Psi}=E_{gs}\ket{\Psi}

from

\hat H'\ket{\Psi}=(\hat H-\hat V+\hat V')\ket{\Psi}=E_{gs}'\ket{\Psi}

we would get V-V'=E-E', which is a contradiction with the assumption that V and V' differ by more than a constant.

Similarly, from the ground state wavefunction \Psi we can compute the charge density n giving rise to the map

D: \Psi \to n

which is also bijective, because to every \Psi we can compute n from (8.9.1.1) and two different \Psi and \Psi' give two different n and n', because different \Psi and \Psi' give

E_{gs}=\braket{\Psi|\hat H|\Psi}<\braket{\Psi'|\hat H|\Psi'}= \braket{\Psi'|\hat H'+\hat V-\hat V'|\Psi'}=E_{gs}'+\int n'({\bf r}) (v({\bf r})-v'({\bf r}))\,\d^3 r

E_{gs}'=\braket{\Psi'|\hat H'|\Psi'}<\braket{\Psi|\hat H'|\Psi}= \braket{\Psi|\hat H+\hat V'-\hat V|\Psi}=E_{gs}+\int n({\bf r}) (v'({\bf r})-v({\bf r}))\,\d^3 r

adding these two inequalities together gives

0<\int n'({\bf r}) (v({\bf r})-v'({\bf r}))\,\d^3 r + \int n({\bf r}) (v'({\bf r})-v({\bf r}))\,\d^3 r= \int (n({\bf r})-n'({\bf r}))(v'({\bf r})-v({\bf r}))\,\d^3 r

which for n=n' gives 0<0, which is nonsense, so n\neq n'.

So we have proved that for a given ground state density n_0({\bf r}) (generated by a potential \hat V_0) it is possible to calculate the corresponding ground state wavefunction \Psi_0({\bf r_1},{\bf r_2},\cdots,{\bf
r_N}), in other words, \Psi_0 is a unique functional of n_0:

\Psi_0=\Psi_0[n_0]

so the ground state energy E_0 is also a functional of n_0

E_0=\braket{\Psi_0[n_0]|\hat T+\hat U+\hat V_0|\Psi_0[n_0]}=E[n_0]

We define an energy functional

(8.9.2.1)E_{v_0}[n]=\braket{\Psi[n]|\hat T+\hat U+\hat V_0|\Psi[n]}= \braket{\Psi[n]|\hat T+\hat U|\Psi[n]}+\int v_0({\bf r})n({\bf r})\d^3r

where \ket{\Psi[n]} is any ground state wavefunction (generated by an arbitrary potential), that is, n is a ground state density belonging to an arbitrary system. E_0 which is generated by the potential V_0 can then be expressed as

E_0=E_{v_0}[n_0]

and for n\neq n_0 we have (from the Ritz principle)

E_0<E_{v_0}[n]

and one has to minimize the functional E_{v_0}[n]:

(8.9.2.2)E_0=\min_n E_{v_0}[n]

The term

\braket{\Psi[n]|\hat T+\hat U|\Psi[n]}\equiv F[n]

in (8.9.2.1) is universal in the sense that it doesn’t depend on \hat
V_0. It can be proven [DFT], that F[n] is a functional of n for degenerated ground states too, so (8.9.2.2) stays true as well.

The ground state densities in (8.9.2.1) and (8.9.2.2) are called pure-state v-representable because they are the densities of (possible degenerate) ground state of the Hamiltonian with some local potential v({\bf
r}). One may ask a question if all possible functions are v-representable (this is called the v-representability problem). The question is relevant, because we need to know which functions to take into account in the minimization process (8.9.2.2). Even though not every function is v-representable [DFT], every density defined on a grid (finite of infinite) which is strictly positive, normalized and consistent with the Pauli principle is ensemble v-representable. Ensemble v-representation is just a simple generalization of the above, for details see [DFT].

The functional E_{v_0}[n] in (8.9.2.2) depends on the particle number N, so in order to get n, we need to solve the variational formulation

{\delta\over\delta n}\left(E_v[n]-\mu(N)\int n(\bf r)\d^3r\right)=0

so

(8.9.2.3){\delta E_v[n]\over\delta n}=\mu(N)

Let the n_N(\bf r) be the solution of (8.9.2.3) with a particle number N and the energy E_N:

E_N=E_v[n_N]

The Lagrangian multiplier \mu is the exact chemical potential of the system

\mu(N)={\partial E_N\over\partial N}

becuase

E_{N+\epsilon}-E_N=E_v[n_{N+\epsilon}]-E_v[n_N] =\int {\delta E_v\over\delta n} (n_{N+\epsilon}-n_N)\d^3r=

=\int \mu(N) (n_{N+\epsilon}-n_N)\d^3r =\mu(N)(N+\epsilon-N)=\mu(N)\epsilon

so

\mu(N)={E_{N+\epsilon}-E_N\over\epsilon} \ \longrightarrow \ {\partial E_N\over\partial N}

8.9.3. The Kohn-Sham Equations

Consider an auxiliary system of N noninteracting electrons (noninteracting gas):

\hat H_s=\hat T+\hat V_s

the Schrödinger equation then becomes:

(-\half\nabla^2+v_s({\bf r}))\psi_i({\bf r}) =\epsilon_i\psi_i({\bf r})

n_s({\bf r})=\sum_i^N|\psi_i({\bf r})|^2

and the total energy is:

E_s[n]=T_s[\{\psi_i[n]\}]+V_s[n]

where

T_s[n]=\braket{\Psi[n]|\hat T|\Psi[n]}= \sum_i\braket{\psi_i|-\half\nabla^2|\psi_i}

V_s[n]=\braket{\Psi[n]|\hat V|\Psi[n]}=\int v_s({\bf r})n({\bf r})\d^3r

So:

E_s[n] = \sum_i\braket{\psi_i|-\half\nabla^2|\psi_i} +
    \int v_s({\bf r})n({\bf r})\d^3r
    =

    = \sum_i\int \psi_i^* \left(-\half\nabla^2\right)\psi_i\,\d^3 r +
        \int v_s({\bf r})\sum_i\psi_i^* \psi_i\, \d^3r
    =

    = \sum_i\int \psi_i^* \left(-\half\nabla^2 + v_s({\bf r})\right)
        \psi_i\,\d^3 r
    =

    = \sum_i \epsilon_i\int \psi_i^* \psi_i\,\d^3 r =

    = \sum_i \epsilon_i

The total energy is the sum of eigenvalues (energies of the individual independent particles) as expected. From the last equation it follows:

T_s[n] = \sum_i\braket{\psi_i|-\half\nabla^2|\psi_i}
    = \sum_i \epsilon_i -\int v_s({\bf r})n({\bf r})\d^3r

In other words, the kinetic energy of the noninteracting particles is equal to the sum of eigenvalues minus the potential energy coming from the total effective potential v_s used to construct the single particle orbitals \psi_i.

From (8.9.2.3) we get

(8.9.3.1)\mu={\delta E_s[n]\over\delta n({\bf r})}= {\delta T_s[n]\over\delta n({\bf r})}+{\delta V_s[n]\over\delta n({\bf r})}= {\delta T_s[n]\over\delta n({\bf r})}+v_s({\bf r})

Solution to this equation gives the density n_s.

Now we want to express the energy in (8.9.1.2) using T_s and E_H for convenience, where E_H is the classical electrostatic interaction energy of the charge distribution \rho({\bf r}), defined using following relations - we start with a Poisson equation in atomic units

\nabla^2 \phi_H({\bf r})=-4\pi \rho({\bf r})

and substitute \rho({\bf r}) = q n({\bf r}), V_H({\bf r}) = q \phi_H({\bf r}) and we use the fact that q^2=1 in atomic units:

\nabla^2 V_H({\bf r})=-4\pi q^2 n({\bf r}) = -4\pi n({\bf r})

or equivalently by expressing V_H using the Green function:

(8.9.3.2)V_H({\bf r})
    = -{1\over 4\pi} \int {-4\pi n({\bf r'})\over|{\bf r}-{\bf r'}|} \d^3r'
    = \int {n({\bf r'})\over|{\bf r}-{\bf r'}|} \d^3r'

and finally E_H is related to V_H using:

V_H({\bf r})={\delta E_H\over\delta n({\bf r})}

so we get:

E_H[n]=\half\int\int {n({\bf r})n({\bf r'})\over|{\bf r}-{\bf r'}|} \d^3r\d^3r'

Using the rules for functional differentiation, we can check that:

V_H({\bf r})
    ={\delta E_H\over\delta n({\bf r})}
    ={\delta \over\delta n({\bf r})}
        \half\int\int {n({\bf r'})n({\bf r''})\over|{\bf r'}-{\bf r''}|}
        \d^3r'\d^3r''
    =

    =\int {n({\bf r'})\over|{\bf r}-{\bf r'}|}
        \d^3r'

Using the above relations, we can see that

E_H[n]=\half\int V_H({\bf r}) n({\bf r}) \d^3r

So from (8.9.2.1) we get

(8.9.3.3)E[n]=(T+U)[n]+V[n]=T_s[n]+E_H[n]+(T-T_s+U-E_H)[n]+V[n]=

=T_s[n]+E_H[n]+E_{xc}[n]+V[n]

The rest of the energy is denoted by E_{xc}=U-E_H+T-T_s and it is called is the exchange and correlation energy functional. From (8.9.2.3)

\mu={\delta E[n]\over\delta n({\bf r})}= {\delta T_s[n]\over\delta n({\bf r})}+ {\delta E_H[n]\over\delta n({\bf r})}+ {\delta E_{xc}[n]\over\delta n({\bf r})}+ {\delta V[n]\over\delta n({\bf r})}

From (8.9.3.2) we have

{\delta E_H\over\delta n({\bf r})}=V_H({\bf r})

from (8.9.1.3) we get

{\delta V[n]\over\delta n({\bf r})}=v({\bf r})

we define

(8.9.3.4){\delta E_{xc}[n]\over\delta n({\bf r})}=V_{xc}({\bf r})

so we arrive at

(8.9.3.5)\mu={\delta E[n]\over\delta n({\bf r})}= {\delta T_s[n]\over\delta n({\bf r})}+V_H({\bf r})+V_{xc}({\bf r})+v({\bf r})

Solution to this equation gives the density n. Comparing (8.9.3.5) to (8.9.3.1) we see that if we choose

(8.9.3.6)v_s\equiv V_H+V_{xc}+v

then n_s({\bf r})\equiv n({\bf r}). So we solve the Kohn-Sham equations of this auxiliary non-interacting system

(8.9.3.7)(-\half\nabla^2+v_s({\bf r}))\psi_i({\bf r}) \equiv(-\half\nabla^2+V_H({\bf
r})+V_{xc}({\bf r})+v({\bf r}))\psi_i({\bf r}) =\epsilon_i\psi_i({\bf r})

which yield the orbitals \psi_i that reproduce the density n({\bf r}) of the original interacting system

(8.9.3.8)n({\bf r})\equiv n_s({\bf r})=\sum_i^N|\psi_i({\bf r})|^2

The sum is taken over the lowest N energies. Some of the \psi_i can be degenerated, but it doesn’t matter - the index i counts every eigenfunction including all the degenerated. In plain words, the trick is in realizing, that the ground state energy can be found by minimizing the energy functional (8.9.2.1) and in rewriting this functional into the form (8.9.3.3), which shows that the interacting system can be treated as a noninteracting one with a special potential.

8.9.4. The XC Term

The exchange and correlation functional

E_{xc}[n]=(T+U)[n]-E_H[n]-T_S[n]

can always be written in the form

E_{xc}[n]=\int n({\bf r}')\epsilon_{xc}({\bf r}';n)\d^3r'

where the \epsilon_{xc}({\bf r}';n) is called the XC energy density. The XC potential is defined as:

V_{xc}({\bf r};n) = {\delta E_{xc}[n]\over\delta n({\bf r})}
    = \epsilon_{xc}({\bf r};n)+ \int n({\bf r}')
        {\delta \epsilon_{xc}({\bf r}';n)\over\delta n({\bf r})}\d^3r'

8.9.5. Total Energy

We already derived all the necessary things above, so we just summarize it here. The total energy is given by:

E[n]=(T+U)[n]+V[n]=T_s[n]+E_H[n]+(T-T_s+U-E_H)[n]+V[n]=

=T_s[n]+E_H[n]+E_{xc}[n]+V[n]

where

T_s[n] = \sum_i \epsilon_i -\int v_s({\bf r})n({\bf r})\d^3r

E_H[n] = \half\int V_H({\bf r}) n({\bf r}) \d^3r

E_{xc}[n]=\int \epsilon_{xc}({\bf r};n) n({\bf r}) \d^3r

V[n]=\int v({\bf r}) n({\bf r}) \d^3r

This is the correct, quadratically convergent expression for the total energy. We use the whole input potential V_{in}\equiv v_s and its associated eigenvalues \epsilon_i to calculate the kinetic energy T_s[n], this follows from the derivation of the expression for T_s[n]. Then we use the calculated charge density to express E_H[n], E_{xc}[n] and V[n].

If one is not careful about the potential associated with the eigenvalues, i.e., confusing V_{in} with V_{out}, one gets a slowly converging formula for the total energy. By expanding v_s using (8.9.3.6):

\int v_s n({\bf r})\d^3 r
    = \int (V_H + V_{xc} + v) n({\bf r})\d^3 r
    = 2 \half\int V_H n({\bf r})\d^3 r
    + \int V_{xc} n({\bf r})\d^3 r
    + \int v n({\bf r})\d^3 r =

    = 2 E_H[n] + \int V_{xc} n({\bf r})\d^3 r + V[n]

So T_s is equal to:

T_s[n] = \sum_i \epsilon_i -\int v_s({\bf r})n({\bf r})\d^3r =

    = \sum_i \epsilon_i - 2 E_H[n] - \int V_{xc} n({\bf r})\d^3 r - V[n]

And then the slowly converging form of total energy is:

E[n] = T_s[n]+E_H[n]+E_{xc}[n]+V[n]
    = \sum_i \epsilon_i - 2 E_H[n] - \int V_{xc} n({\bf r})\d^3 r - V[n]
        +E_H[n]+E_{xc}[n]+V[n] =

    = \sum_i \epsilon_i - E_H[n] + E_{xc}[n]
        -\int V_{xc}({\bf r};n) n({\bf r}) \d^3r

The reason it is slowly converging is because the new formula for kinetic energy is mixing V_{in} with V_{out}, so it is not as precise (see above) and converges much slower with SCF iterations. Once self-consistency has been achieved (i.e. V_{in}=V_{out}), the two expressions for total energy are equivalent.

8.9.6. XC Approximations

All the expressions above are exact (no approximation has been made so far). Unfortunately, no one knows \epsilon_{xc}({\bf r}';n) exactly (yet). As such, various approximations for it exist.

LDA

The most simple approximation is the local density approximation (LDA), for which the xc energy density \epsilon_{xc} at \bf r is taken as that of a homogeneous electron gas (the nuclei are replaced by a uniform positively charged background, density n=\rm const) with the same local density:

\epsilon_{xc}({\bf r};n)\approx\epsilon_{xc}^{LD}(n({\bf r}))

The xc potential V_{xc} defined by (8.9.3.4) is then

V_{xc}({\bf r};n)={\delta E_{xc}[n]\over\delta n({\bf r})}= \epsilon_{xc}({\bf r};n)+ \int n({\bf r}'){\delta \epsilon_{xc}({\bf r}';n)\over\delta n({\bf r})}\d^3r'

which in the LDA becomes

(8.9.6.1)V_{xc}({\bf r};n) =\epsilon_{xc}^{LD}(n)+n{\d \epsilon_{xc}^{LD}(n)\over \d n}= {\d \over \d n}\left(n\epsilon_{xc}^{LD}(n)\right)= V_{xc}^{LD}(n)

The xc energy density \epsilon_{xc}^{LD} of the homogeneous gas can be computed exactly:

\epsilon_{xc}^{LD}(n)=\epsilon_x^{LD}(n)+\epsilon_c^{LD}(n)

where the \epsilon_x^{LD} is the electron gas exchange term given by

\epsilon_x^{LD}(n)=-{3\over4\pi}(3\pi^2 n)^{1\over3}

the rest of \epsilon_{xc}^{LD} is hidden in \epsilon_c^{LD}(n) for which there doesn’t exist an analytic formula, but the correlation energies are known exactly from quantum Monte Carlo (QMC) calculations by Ceperley and Alder [pickett]. The energies were fitted by Vosko, Wilkes and Nussair (VWN) with \epsilon_c^{LD}(n) and they got accurate results with errors less than 0.05\rm\,mRy in \epsilon_c^{LD}, which means that \epsilon_c^{LD}(n) is virtually known exactly. VWN result:

\epsilon_c^{LD}(n)\approx {A\over2}\left\{ \ln\left(y^2\over Y(y)\right)+{2b\over Q}\arctan\left(Q\over 2y+b\right)+ \right.

\left. -{by_0\over Y(y_0)}\left[\ln\left((y-y_0)^2\over Y(y)\right) +{2(b+2y_0)\over Q}\arctan\left(Q\over 2y+b\right) \right] \right\}

where y=\sqrt{r_s}, Y(y)=y^2+by+c, Q=\sqrt{4c-b^2}, y_0=-0.10498, b=3.72744, c=12.9352 (note that the value of c is wrong in [pickett]), A=0.0621814 and r_s is the electron gas parameter, which gives the mean distance between electrons (in atomic units):

r_s=\left(3\over4\pi n\right)^{1\over3}

The xc potential is then computed from (8.9.6.1):

V_{xc}^{LD}=V_x^{LD}+V_c^{LD}

V_x^{LD}=-{1\over\pi}(3\pi^2 n)^{1\over3} = {4\over 3}\epsilon_x^{LD}

V_c^{LD}={A\over2}\left\{ \ln\left(y^2\over Y(y)\right)+{2b\over Q}\arctan\left(Q\over 2y+b\right)+ \right.

\left. -{by_0\over Y(y_0)}\left[\ln\left((y-y_0)^2\over Y(y)\right) +{2(b+2y_0)\over Q}\arctan\left(Q\over 2y+b\right) \right] \right\}+

-{A\over6}{c(y-y_0)-by_0y\over (y-y_0)Y(y)}

Some people also use Perdew and Zunger formulas, but they give essentially the same results. The LDA, although very simple, is surprisingly successful. More sophisticated approximations exist, for example the generalized gradient approximation (GGA), which sometimes gives better results than the LDA, but is not perfect either. Other options include orbital-dependent (implicit) density functionals or a linear response type functionals, but this topic is still evolving. The conclusion is, that the LDA is a good approximation to start with, and only when we are not satisfied, we will have to try some more accurate and modern approximation.

RLDA

Relativistic corrections to the energy-density functional (RLDA) were proposed by MacDonald and Vosko:

\epsilon_x^{RLD}(n) = \epsilon_x^{LD}(n)R

R = 1-{3\over2}\left(\beta\mu-\ln(\beta+\mu)\over\beta^2\right)^2
= 1-{3\over2} A^2

where

\mu=\sqrt{1+\beta^2}

\beta={(3\pi^2n)^{1\over3}\over c} = -{4\pi\over 3c} \epsilon_x^{LD}

A = {\beta\mu-\log(\beta+\mu)\over\beta^2}

We now calculate V_x^{RLD}:

(8.9.6.2)V_x^{RLD}
=\epsilon_x^{LD}R+n{\d \epsilon_x^{LD}R\over\d n}=

=\epsilon_x^{LD}R+n{\d \epsilon_x^{LD}\over\d n}R
    +n\epsilon_x^{LD}{\d R\over\d n}
    =

=\epsilon_x^{LD}R+n{\d \epsilon_x^{LD}\over\d n}R
    +n\epsilon_x^{LD}{\d R\over\d \beta}{\d\beta\over\d n}

where the derivative {\d\beta\over\d n} can be evaluated as follows:

{\d \beta\over\d n}
    = {\d \over \d n}{(3\pi^2n)^{1\over3}\over c}
    = {1 \over 3n}{(3\pi^2n)^{1\over3}\over c}
    = {\beta\over 3n}

And {\d \epsilon_x^{LD}\over\d n} in exactly the same manner:

{\d \epsilon_x^{LD}\over\d n} = ... = {\epsilon_x^{LD}\over 3n}

So we can write

V_x^{RLD}
=\epsilon_x^{LD}R+n{\d \epsilon_x^{LD}\over\d n}R
    +n\epsilon_x^{LD}{\d R\over\d \beta}{\d\beta\over\d n}
=

=\epsilon_x^{LD}R+n{\epsilon_x^{LD}\over 3n}R
    +n\epsilon_x^{LD}{\d R\over\d \beta}{\beta\over 3n}
=

=
{4\over3}\epsilon_x^{LD}R+{1\over3}\epsilon_x^{LD}{\d R\over\d\beta}\beta
=

=
{4\over3}\epsilon_x^{LD}\left(R+{1\over4}\beta{\d R\over\d\beta}\right)
=

=
V_x^{LD}\left(R+{1\over4}\beta{\d R\over\d\beta}\right)

where

{\d R\over\d \beta}= -{3\over 2} 2 A A' = -3AA'
=

=-6 A \left({1\over\mu} - {A\over\beta}\right)

where we used the derivative of A(\beta), which after a tedious, but straightforward differentiation is:

A'(\beta) = \cdots = 2\left({1\over\mu} - {A\over\beta}\right)

Plugging this back in, we get:

V_x^{RLD} = V_x^{LD}\left(R+{1\over4}\beta{\d R\over\d\beta}\right)
=

=
V_x^{LD}\left(1-{3\over 2}A^2 +{1\over4}\beta(-6A)
\left({1\over\mu} - {A\over\beta}\right)
\right)
=

=
V_x^{LD}\left(1-{3\over 2}A^2 + {6\over4} A^2-{6\over4}\beta {A\over\mu}
\right)
=

=
V_x^{LD}\left(1-{3\over2}{\beta\over\mu} A
\right)
=

=
V_x^{LD}\left(1-{3\over2}{\beta\over\mu} \left(
    \beta\mu - \log(\beta+\mu) \over \beta^2\right)
\right)
=

=
V_x^{LD}\left(1-{3\over2} \left(
    \beta\mu - \log(\beta+\mu) \over \beta\mu\right)
\right)
=

= V_x^{LD}\left({3\log(\beta+\mu) \over 2\beta\mu}-\half\right)

For c\to\infty we get \beta\to0, R\to1 and V_x^{RLD}\to
{4\over3}\epsilon_x^{LD}=V_x^{LD} as expected, because

\lim_{\beta\to0}{\beta\sqrt{1+\beta^2}-\ln(\beta+\sqrt{1+\beta^2})\over \beta^2} = 0

Code:

>>> from sympy import limit, var, sqrt, log
>>> var("beta")
beta
>>> limit((beta*sqrt(1+beta**2) - log(beta+sqrt(1+beta**2)))/beta**2, beta, 0)
0

8.9.7. Radial DFT Problem

Kohn-Sham Equations

For spherically symmetric potentials, we write all eigenfunctions as:

\psi_{nlm} = R_{nl} Y_{lm}

and we need to solve the following Kohn-Sham equations:

-{1\over2}R_{nl}'' - {1\over r}R_{nl}' +
    \left(V + {l(l+1)R\over2 r^2}\right)R_{nl} = \epsilon_{nl} R_{nl}

With normalization:

\int R_{nl}^2 \,r^2 \,\d r = 1

For Schroedinger equation, the charge density is calculated by adding all “(n, l, m)” states together, counting each one twice (for spin up and spin down):

n({\bf r}) = \sum_{nlm}2|\psi_{nlm}|^2
    = \sum_{nlm}R_{nl}^2 2|Y_{lm}|^2
    = \sum_{nl}R_{nl}^2 2\sum_m|Y_{lm}|^2
    = {1\over 4\pi}\sum_{nl}f_{nl} R_{nl}^2

where we have introduced the occupation numbers f_{nl} by

f_{nl} = 4\pi\,2\sum_m |Y_{lm}|^2

Normalization of the charge density is:

Z = \int n({\bf r}) \d^3 x
    = \int n(r) \, r^2\d\Omega \d r
    = 4\pi\int n(r) \, r^2 \d r
    =

    = 4\pi\int {1\over 4\pi} \sum_{nl} f_{nl} R_{nl}^2\, r^2\d\Omega \d r
    =

    = \sum_{nl} f_{nl}\int R_{nl}^2\, r^2 \d r
    =

    = \sum_{nl} f_{nl}

So we can see, that it must hold:

\sum_{nl} f_{nl} = Z

where Z is the atomic number (number of electrons), and as such, f_{nl} are indeed the occupation numbers (integers). The factor 4\pi is explicitly factored out, as it cancels with the spherical harmonics: assuming all m states are occupied, this can be simplified to:

f_{nl} = 4\pi\,2\sum_m |Y_{lm}|^2
    = 4\pi\,2{2l+1\over 4\pi}
    = 2(2l+1)

We can also use this machinery to prescribe “chemical occupation numbers”, that don’t necessarily correspond to the DFT ground state. For example for U atom we get:

f_{1l} = 2 (2l+1)

f_{2l} = 2 (2l+1)

f_{3l} = 2 (2l+1)

f_{4l} = 2 (2l+1)

f_{5l} = 2 (2l+1)\quad\quad l\le2

f_{53} = 3

f_{60} = 2

f_{61} = 6

f_{62} = 1

f_{70} = 2

By summing all these f_{nl}, we get 92 as expected:

\sum_{nl} f_{nl} = 2 + (2+6) + (2+6+10) + (2+6+10+14) + (2+6+10) +

    + 3 + 2 + 6 + 1 + 2 = 92

But this isn’t the DFT ground state, because some KS energies are skipped, for example there is only one state for n=6, l=2, but there are nine more states with the same energy — instead two more states are occupied in n=7, l=0, but those have higher energy. So this corresponds to excited DFT state, strictly speaking not physically valid in the DFT formalism, but in practice this approach is often used. One can also prescribe fractional occupation numbers (in the Dirac case).

Poisson Equation

Poisson equation becomes:

V_H''(r) + {2\over r} V_H'(r) = -4\pi n(r)

Total Energy

The total energy is given by:

E[n]= T_s[n]+E_H[n]+E_{xc}[n]+V[n]

where

T_s[n] = \sum_{nl} f_{nl}\epsilon_{nl}
    -\int (V_H(r) + V_{xc}(r) + v(r))_{in} n(r) \d^3 r
    =

    = \sum_{nl} f_{nl}\epsilon_{nl}
        -\int \left(V_H(r) + V_{xc}(r) -{Z\over r}\right)_{in} n(r) \d^3 r

E_H[n] = \half\int V_H(r) n(r) \d^3r

E_{xc}[n]=\int \epsilon_{xc}(r;n) n(r) \d^3r

V[n]=\int v(r) n(r) \d^3r = -\int {Z\over r} n(r) \d^3r

doing the integrals a bit we get:

T_s[n] = \sum_{nl} f_{nl}\epsilon_{nl}
        -4\pi\int \left(V_H(r) + V_{xc}(r) -{Z\over r}\right)_{in} n(r)
            r^2\,\d r

E_H[n] = 2\pi\int V_H(r) n(r)r^2\, \d r

E_{xc}[n]=4\pi\int \epsilon_{xc}(r;n) n(r)r^2\, \d r

V[n]=-4\pi \int {Z\over r} n(r)r^2\, \d r
    =-4\pi Z \int n(r)r\, \d r

We can also express everything using the charge density \rho(r) = -n(r):

T_s[n] = \sum_{nl} f_{nl}\epsilon_{nl}
        +4\pi\int \left(V_H(r) + V_{xc}(r) -{Z\over r}\right)_{in} \rho(r)
            r^2\,\d r

E_H[n] = -2\pi\int V_H(r) \rho(r)r^2\, \d r

E_{xc}[n]=-4\pi\int \epsilon_{xc}(r;n) \rho(r)r^2\, \d r

V[n]=4\pi \int {Z\over r} \rho(r)r^2\, \d r
    =4\pi Z \int \rho(r)r\, \d r

8.9.8. DFT As a Nonlinear Problem

The task is to find such a charge density n, so that all the equations below hold (e.g. are self-consistent):

V = -{Z\over r} + V_H + V_{xc}

\left(-\nabla^2+V\right)\phi_m = \epsilon_m\phi_m,\quad\quad m = 1, 2, \dots, 4

n = \sum_{m=1}^4 \phi_m^2

V_{xc} = f(n)

\nabla^2 V_H = -4\pi n

This is a standard nonlinear problem, except that the Jacobian is dense, as shown below.

Reformulation

Let’s write everything in terms of \phi_m(x) explicitly:

n(x) = \sum_{m=1}^4 \phi_m^2(x)

V_{xc}(x) = f(n(x)) = f\left( \sum_{m=1}^4\phi_m^2(x) \right)

V_H(x) = \int_\Omega {n(x')\over|x' - x|}\d x'=
\int_\Omega {
\sum_{m=1}^4 \phi_m^2(x')
\over|x' - x|}\d x'

V(x) = -{Z\over r} + V_H(x) + V_{xc}(x)=

=-{Z\over r}+
\int_\Omega {
\sum_{m=1}^4 \phi_m^2(x')
\over|x' - x|}\d x'
+f\left( \sum_{m=1}^4\phi_m^2(x) \right)

Now we can write everything as just one (nonlinear) equation:

\left(-\nabla^2
-{Z\over r}+
\int_\Omega {
\sum_{m=1}^4 \phi_m^2(x')
\over|x' - x|}\d x'
+f\left( \sum_{m=1}^4\phi_m^2(x) \right)
\right)\phi_n = \epsilon_n\phi_n,\quad\quad n = 1, 2, \dots, 4

FE Discretization

The correspondig discrete problem has the form

\int_\Omega \nabla\phi_n(x)\cdot\nabla v_i(x)+\left[
-{Z\over r}+
\int_\Omega {
\sum_{m=1}^4 \phi_m^2(x')
\over|x' - x|}\d x'
+f\left( \sum_{m=1}^4\phi_m^2(x) \right)
\right]
\phi_n(x) v_i(x)  \d x=

=\int_\Omega
\epsilon_n\phi_n(x) v_i(x) \d x,\quad\quad n = 1, 2, \dots, 4;\quad
i = 1, 2, \dots, N

where

\phi_n = \phi_n({\bf Y}^{(n)}) = \sum_{j=1}^N y_j^{(n)} v_j(x)

Here {\bf Y}^{(n)} = (y_1^{(n)}, y_2^{(n)}, \dots, y_N^{(n)})^T is the vector of unknown coefficients for the n-th wavefunction \phi_n(x). Our equation can then be written in the compact form

{\bf F}({\bf Y}^{(n)}) = {\bf 0},\quad\quad n = 1, 2, \dots, 4

where {\bf F} = (F_1, F_2, \dots, F_N)^T with

F_i({\bf Y}^{(n)}) =
\int_\Omega \nabla\phi_n(x)\cdot\nabla v_i(x)+\left[
-{Z\over r}+
\int_\Omega {
\sum_{m=1}^4 \phi_m^2(x')
\over|x' - x|}\d x'
+f\left( \sum_{m=1}^4\phi_m^2(x) \right)
\right]
\phi_n(x) v_i(x)  \d x-

-\int_\Omega
\epsilon_n\phi_n(x) v_i(x) \d x

Jacobian

The Jacobi matrix has the elements:

J_{ik} = {\partial F_i\over\partial y_k^{(s)}}

The only possible dense term is:

{\partial\over\partial y_k^{(s)}}\int_\Omega \int_\Omega {
\sum_{m=1}^4 \phi_m^2(x')
\over|x' - x|}\d x'\,\phi_n(x) v_i(x) \d x =

=
{\partial\over\partial y_k^{(s)}}\int_\Omega \int_\Omega {
\sum_{m=1}^4 \left(\sum_{j=1}^N y_j^{(m)} v_j(x')\right)^2
\over|x' - x|}\d x'\, \left(\sum_{j=1}^N y_j^{(n)} v_j(x)\right)  v_i(x) \d x =

=
\int_\Omega \int_\Omega {
2 \left(\sum_{j=1}^N y_j^{(s)} v_j(x')\right)v_k(x')
\over|x' - x|}\d x'\, \left(\sum_{j=1}^N y_j^{(n)} v_j(x)\right)  v_i(x) \d x +

+
\int_\Omega \int_\Omega {
\sum_{m=1}^4 \left(\sum_{j=1}^N y_j^{(m)} v_j(x')\right)^2
\over|x' - x|}\d x'\, \delta_{ns}v_k(x)  v_i(x) \d x

Now we can see that we have in there the following term:

\int_\Omega \int_\Omega {v_k(x') v_i(x)\over |x'-x|}\d x'\d x

which is dense in (ki), as can be easily seen be fixing i and writing

\int_\Omega \int_\Omega {v_k(x')\over |x'-x|}\d x' v_i(x)\d x

so for each k there is some contribution from the integral \int_\Omega
{v_k(x')\over |x'-x|}\d x' for such x where v_i(x) is nonzero, thus making the Jacobian J_{ik} dense.

8.9.9. Thomas-Fermi-Dirac Theory

There are two ways to derive equations for Thomas-Fermi-Dirac theory. One way is to start from grand potential and derive all equations from it. The other way is to start with low level equations and build our way up. Will start with the former.

Top Down Approach

We start with a grand potential for fermions:

\Omega[\beta, \mu]
= -\sum_i {1\over\beta}
    \log\left(\sum_{N=0}^1 e^{-\beta\left(N\epsilon_i - N\mu\right)}\right)
        =

= -\sum_i {1\over\beta}
    \log\left(1 + e^{-\beta\left(\epsilon_i - \mu\right)}\right)
        =

= -{1\over\beta}
    \int \int {2\d^3 x \d^3 p \over (2\pi)^3} \log\left(1 +
        e^{-\beta\left({p^2\over 2} + V({\bf x}) - \mu\right)}\right)
        -E_{ee} - {1\over3}E_{xc} =

= -{2\over\beta}
    \int \d^3 x \int_0^\infty{ 4\pi p^2 \d p \over (2\pi)^3} \log\left(1 +
        e^{-\beta\left({p^2\over 2} + V({\bf x}) - \mu\right)}\right)
        -E_{ee} - {1\over3}E_{xc} =

= -{1\over \pi^2 \beta}
    \int \d^3 x \int_0^\infty p^2 \log\left(1 +
        e^{-\beta\left({p^2\over 2} + V({\bf x}) - \mu\right)}\right) \d p
        -E_{ee} - {1\over3}E_{xc} =

= -{2\sqrt2 \over 3 \pi^2 \beta^{5\over2}}
    \int \d^3 x \int_0^\infty {u^{3\over2} \over
        1 + e^{u-\beta\left(\mu-V({\bf x})\right)}} \d u
            -E_{ee} - {1\over3}E_{xc} =

= -{2\sqrt2 \over 3 \pi^2 \beta^{5\over2}}
    \int I_{3\over2}\left(\beta\left(\mu-V({\bf x})\right)\right) \,\d^3 x
    -E_{ee} - {1\over3}E_{xc} =

= -{2\sqrt2 \over 3 \pi^2 \beta^{5\over2}}
    \int I_{3\over2}\left(\beta\left(\mu-V({\bf x})\right)\right) \,\d^3 x
    -E_{ee} - {1\over3}E_{xc}

The potential V({\bf x}) = V_{en}({\bf x}) + V_{ee}({\bf x}) + V_{xc}({\bf x}) is the total potential that the electrons experience (it contains nuclear, Hartree, and XC terms) and E_{ee} is the Hartree energy:

E_{en} = \int n_e({\bf x}) V_{en}({\bf x})\,\d^3 x\,,

E_{ee}
    = \half \int n_e(\mathbf{x}) V_{ee}(\mathbf{x}) \d^3 x\,,

E_{xc} = {3\over 4}\int n_e({\bf x}) V_{xc}({\bf x})\,\d^3 x\,.

For simplicity, we assume here that V_{xc} only contains the exchange of the homogeneous electron gas. For a general XC functional, the relation is nonlinear and one must simply numerically calculate the XC energy density e_{xc}({\bf x}) and calculate the XC energy using:

E_{xc} = \int n_e({\bf x}) e_{xc}({\bf x})\,\d^3 x\,.

In our case here, we have e_{xc} = {3\over4}V_{xc}({\bf x}), which is only true for the exchange in homogeneous electron gas. Otherwise the relation is nonlinear. In the general case, the correction that must be applied is:

E_{en} + E_{ee} + E_{xc} - \int n_e({\bf x}) V({\bf x})\,\d^3 x =

= \int n_e({\bf x})\left(
    (e_{en}-V_{en})
    +(e_{ee}-V_{ee})
    +(e_{xc}-V_{xc})
    \right)\,\d^3 x =

= \int n_e({\bf x})\left(
    0
    +\left(\half V_{ee}-V_{ee}\right)
    +\left({3\over 4}V_{xc}-V_{xc}\right)
    \right)\,\d^3 x =

= \int n_e({\bf x})\left(
    -\half V_{ee}
    -{1\over 4}V_{xc}
    \right)\,\d^3 x =

= -\half\int n_e({\bf x}) V_{ee} \d^3 x
- {1\over 3}{3\over 4}\int n_e({\bf x}) V_{xc} \d^3 x =

= -E_{ee} - {1\over 3} E_{xc}

The density is a functional derivative with respect to \mu:

n_e({\bf x}) = - {\delta \Omega[\beta, \mu] \over \delta \mu}
    = {2\sqrt2 \over 3 \pi^2 \beta^{5\over2}}
        {\partial \over \partial \mu}
            I_{3\over2}\left(\beta\left(\mu-V({\bf x})\right)\right)
    = {2\sqrt2 \over 3 \pi^2 \beta^{5\over2}}
        \beta {3\over 2} I_{1\over2}
            \left(\beta\left(\mu-V({\bf x})\right)\right) =

    = {\sqrt2 \over \pi^2 \beta^{3\over2}} I_{1\over2}
            \left(\beta\left(\mu-V({\bf x})\right)\right)

By defining the function \Phi(n_e({\bf x})):

\Phi(n_e({\bf x})) = \beta\left(\mu-V({\bf x})\right)
    = I_{1\over2}^{-1}\left(
            {\pi^2 \beta^{3\over2} \over \sqrt 2} n_e({\bf x})
        \right)

we can express the grand potential using n_e as follows:

\Omega[\beta, n_e]
    = -{2\sqrt2 \over 3 \pi^2 \beta^{5\over2}}
        \int I_{3\over2}(\Phi(n_e({\bf x}))) \, \d^3 x
      - \half \int n_e(\mathbf{x}) V_{ee}(\mathbf{x}) \d^3
      - {1\over 4} \int n_e(\mathbf{x}) V_{xc}(\mathbf{x}) \d^3 x\,.

Now we can calculate the free energy:

F_e[\beta, n_e] = \Omega[\beta, n_e] + \mu N
    = \Omega[\beta, n_e] + \mu \int n_e({\bf x}) \,\d^3 x =

    = \int \left(-{2\sqrt2 \over 3 \pi^2 \beta^{5\over2}}
          I_{3\over2}(\Phi(n_e({\bf x})))
        + \mu n_e({\bf x})
        - n_e(\mathbf{x}) \left(
            \half V_{ee}(\mathbf{x})
            +{1\over 4} V_{xc}(\mathbf{x})
            \right)
            \right)\d^3 x =

    = \int \left(-{2\sqrt2 \over 3 \pi^2 \beta^{5\over2}}
          I_{3\over2}(\Phi(n_e({\bf x})))
        + {1\over \beta} n_e({\bf x}) \Phi(n_e({\bf x}))
            + n_e({\bf x}) V({\bf x})
            - n_e(\mathbf{x}) \left(
                \half V_{ee}(\mathbf{x})
                +{1\over 4} V_{xc}(\mathbf{x})
                \right)
            \right)\d^3 x =

    = \int \left(-{2\sqrt2 \over 3 \pi^2 \beta^{5\over2}}
          I_{3\over2}(\Phi(n_e({\bf x})))
        + {1\over \beta} n_e({\bf x}) \Phi(n_e({\bf x}))
            + n_e({\bf x})\left( V_{en}({\bf x})
              +\half V_{ee}(\mathbf{x})
              + {3\over 4} V_{xc}(\mathbf{x}) \right)
            \right)\d^3 x\,,

where we used the fact that \mu = {1\over \beta} \Phi(n_e({\bf x})) + V({\bf
x}), i.e. the left hand side \mu is a constant, thus the sum of the terms on the right hand side is also constant (even though the individual terms are not).

We can calculate the entropy S=-\left(\partial\Omega\over\partial T\right)_{V,\mu} as follows:

TS
    =-T \left(\partial\Omega\over\partial T\right)_{V,\mu} =

    =\beta \left(\partial\Omega\over\partial \beta\right)_{V,\mu} =

    =\beta {\partial\over\partial \beta}\left(
        -{2\sqrt2 \over 3 \pi^2 \beta^{5\over2}}
        \int I_{3\over2}(\Phi(n_e({\bf x}))) \, \d^3 x
        - E_{ee} - {1\over3}E_{xc}
    \right) =

    =\beta {\partial\over\partial \beta}\left(
        -{2\sqrt2 \over 3 \pi^2 \beta^{5\over2}}
        \int I_{3\over2}(\Phi(n_e({\bf x}))) \, \d^3 x
    \right) =

    =\beta \left(
        {5\over2}{2\sqrt2 \over 3 \pi^2 \beta^{7\over2}}
        \int I_{3\over2}(\Phi(n_e({\bf x}))) \, \d^3 x
        -{2\sqrt2 \over 3 \pi^2 \beta^{5\over2}}
        \int {3\over2} I_{1\over2}(\Phi(n_e({\bf x})))
            {\partial\Phi(n_e({\bf x}))\over\partial\beta}
        \, \d^3 x
    \right) =

    =\beta \left(
        {5\over2}{2\sqrt2 \over 3 \pi^2 \beta^{7\over2}}
        \int I_{3\over2}(\Phi(n_e({\bf x}))) \, \d^3 x
        -{2\sqrt2 \over 3 \pi^2 \beta^{5\over2}}
        \int {3\over2} I_{1\over2}(\Phi(n_e({\bf x})))
            (\mu-V({\bf x}))
        \, \d^3 x
    \right) =

    = {5\over2}{2\sqrt2 \over 3 \pi^2 \beta^{5\over2}}
        \int I_{3\over2}(\Phi(n_e({\bf x}))) \, \d^3 x
        - \int n_e({\bf x}) (\mu-V({\bf x})) \, \d^3 x =

    = {5\over3}{\sqrt2 \over \pi^2 \beta^{5\over2}}
        \int I_{3\over2}(\Phi(n_e({\bf x}))) \, \d^3 x
        -\mu N + E_{en}+2E_{ee} + {4\over 3}E_{xc}

The total energy is then equal to:

E = \Omega + \mu N + TS =

    = \left(-{2\sqrt2 \over 3 \pi^2 \beta^{5\over2}}
        \int I_{3\over2}(\Phi(n_e({\bf x}))) \, \d^3 x
        - E_{ee} - {1\over3}E_{xc}\right)
        + \mu N
        +{5\over3}{\sqrt2 \over \pi^2 \beta^{5\over2}}
        \int I_{3\over2}(\Phi(n_e({\bf x}))) \, \d^3 x
        -\mu N + E_{en}+2E_{ee} + {4\over 3}E_{xc} =

    = {\sqrt2 \over \pi^2 \beta^{5\over2}}
        \int I_{3\over2}(\Phi(n_e({\bf x}))) \, \d^3 x
        + E_{en} + E_{ee} + E_{xc}

From which we can see that the kinetic energy E_{kin} is equal to:

E_{kin} = E - (E_{en} + E_{ee} + E_{xc}) =

    = {\sqrt2 \over \pi^2 \beta^{5\over2}}
        \int I_{3\over2}(\Phi(n_e({\bf x}))) \, \d^3 x

The relation between the total energy and free energy can be also written as:

E = F + TS = F +
    \beta \left(\partial\Omega\over\partial \beta\right)_{V,\mu} =

    = F + \beta \left(\partial F\over\partial \beta\right)_{V,\mu}
    = \left(\partial (\beta F)\over\partial \beta\right)_{V,\mu}

But it gives the same result as we obtained above.

To determine the kinetic part of the free energy, we set all potentials equal to zero (V({\bf x}) = V_{en}({\bf x}) = V_{ee}({\bf x}) = V_{xc}({\bf x}) =
0) and obtain:

F_{kin}[\beta, n_e]
    = \int \left(-{2\sqrt2 \over 3 \pi^2 \beta^{5\over2}}
          I_{3\over2}(\Phi(n_e({\bf x})))
        + {1\over \beta} n_e({\bf x}) \Phi(n_e({\bf x}))
            \right)\d^3 x\,.

If the potentials are zero, then the pressure can be calculated from:

P = -{1\over V}\Omega[\beta, n_e]
    = {2\sqrt2 \over 3 \pi^2 \beta^{5\over2}V}
        \int I_{3\over2}(\Phi(n_e({\bf x}))) \,\d^3 x =

= {2\sqrt2 \over 3 \pi^2 \beta^{5\over2}V}
        \int I_{3\over2}(\beta\mu) \,\d^3 x
= {2\sqrt2 \over 3 \pi^2 \beta^{5\over2}} I_{3\over2}(\beta\mu) \,.

If the potentials are not zero, then one can calculate the pressure using:

P = - \left(\partial\Omega\over\partial V\right)_{\mu,T}
    = - \left(\partial F\over\partial V\right)_{T,N} =

    = - {\partial \over\partial V} \int f \d^3 x =

    = - \left[f+e_{ee}\right]_b
        - \int {\partial f\over\partial n_e}
          {\partial n_e\over\partial V}
          \d^3 x =

    = - \left[f+e_{ee}\right]_b
        - \mu \int {\partial n_e\over\partial V} \d^3 x =

    = - \left[f+e_{ee}\right]_b
        + \mu [n_e]_b =

    = \left[(-f)-e_{ee}+\mu n_e \right]_b =

    = \left[\left({2\over3}e_{kin} + e_{ee} + {1\over3}e_{xc}-\mu n_e\right)
        -e_{ee}+\mu n_e \right]_b =

    = \left[{2\over3}e_{kin} + {1\over3}e_{xc}\right]_b =

    = {1\over 3V} \int_b \left( {2\over3}e_{kin} + {1\over3}e_{xc}
        \right) {\bf x}\cdot{\bf n}\,\d S =

    = {1\over 3V} \int \left( {2\over3}e_{kin} + {1\over3}e_{xc}
        \right) \nabla\cdot{\bf x}\,\d^3 x
        +
    {1\over 3V} \int {\bf x}\cdot\nabla \left(
        {2\over3}e_{kin} + {1\over3}e_{xc}
        \right) \,\d^3 x =

    = {1\over 3V} (2E_{kin} + E_{xc})
        +
    {1\over 3V} \int {\bf x}\cdot \left(
        -n_e({\bf x})\nabla V({\bf x})
        + \nabla{1\over3}e_{xc}
        \right) \,\d^3 x =

    = {1\over 3V} (2E_{kin} + E_{xc})
        +
    {1\over 3V} (E_{en}+E_{ee}) =

    = {1\over 3V}(2E_{kin} + E_{en} + E_{ee} + E_{xc})

Summary:

\Omega = -{2\over 3} E_{kin} - E_{ee} - {1\over3}E_{xc}

F_e = \Omega + \mu N = -{2\over 3} E_{kin} - E_{ee} - {1\over3}E_{xc}
    + \mu N

TS = {5\over3} E_{kin} + E_{en} + 2 E_{ee} + {4\over3}E_{xc} - \mu N

E = F + TS = \Omega + \mu N + TS = E_{kin} + E_{en} + E_{ee} + E_{xc}\,,

where:

E_{kin} = {\sqrt2 \over \pi^2 \beta^{5\over2}}
        \int I_{3\over2}(\Phi(n_e({\bf x}))) \, \d^3 x

E_{en} = \int n_e({\bf x}) V_{en}({\bf x})\, \d^3 x

E_{ee} = \half \int n_e({\bf x}) V_{ee}({\bf x})\, \d^3 x

E_{xc} = {3\over4}\int n_e({\bf x}) V_{xc}({\bf x})\, \d^3 x

n_e({\bf x}) = {\sqrt2 \over \pi^2 \beta^{3\over2}}
        I_{1\over2}\left( \beta\left(\mu-V({\bf x})\right) \right)

\Phi(n_e({\bf x})) = \beta\left(\mu-V({\bf x})\right)
    = I_{1\over2}^{-1}\left(
            {\pi^2 \beta^{3\over2} \over \sqrt 2} n_e({\bf x})
        \right)

N = \int n_e({\bf x})\, \d^3 x

\mu = {1\over \beta} \Phi(n_e({\bf x})) + V({\bf x})

and \mu N is calculated as follows:

\mu N = \int \mu n_e({\bf x})\, \d^3 x =

    = {1\over \beta} \int \Phi(n_e({\bf x})) n_e({\bf x})\, \d^3 x
        + \int V({\bf x}) n_e({\bf x})\, \d^3 x =

    = {1\over \beta} \int \Phi(n_e({\bf x})) n_e({\bf x})\, \d^3 x
        + E_{en} + 2 E_{ee} + {4\over3} E_{xc} \,.

So F_e can also be expressed as:

F_e = -{2\over 3} E_{kin} - E_{ee} - {1\over3}E_{xc} + \mu N =

    = -{2\over 3} E_{kin}
       + {1\over \beta} \int \Phi(n_e({\bf x})) n_e({\bf x})\, \d^3 x
        + E_{en} + E_{ee} + E_{xc} \,.

Bottom Up Approach

The other way to derive these equations is to use the following considerations. The number of states in a box of side L is given by:

N = \int {\d^3 p \over h^3} 2 L^3
  = \int {\d^3 p \over (2\pi\hbar)^3} 2 L^3
  = \int {\d^3 p \over (2\pi)^3} 2 L^3
  = \int_0^{p_f} {4\pi p^2 \d p\over (2\pi)^3} 2 L^3
  = {L^3\over\pi^2} \int_0^{p_f} p^2 \d p

We use atomic units, so \hbar=1. The electronic particle density is:

(8.9.9.1)n_e({\bf x}) = {N \over L^3}
  = {1\over\pi^2} \int_0^{p_f} p^2 \d p
  = {p_f^3 \over 3\pi^2}
  = {\left[2(E_f - V({\bf x}))\right]^{3\over2} \over 3\pi^2}

where we used the relation for Fermi energy E_f = {p_f^2\over 2} + V({\bf
x}). The potential V({\bf x}) is the total potential that the electrons experience (it contains Hartree, nuclear and XC terms). At finite temperature T we need to use the Fermi distribution and this generalizes to:

n_e({\bf x})
  = {1\over\pi^2} \int_0^{\infty} {p^2 \d p \over
        e^{\beta(E-\mu)} + 1}

Now we use the relation E = {p^2\over 2} + V({\bf x}) and substitutions \epsilon={p^2\over 2}, y = \beta \epsilon to rewrite this using the Fermi-Dirac Integral:

n_e({\bf x})
  = {1\over\pi^2} \int_0^{\infty} {p^2 \d p \over
        e^{\beta(E-\mu)} + 1}
  = {1\over\pi^2} \int_0^{\infty} {p^2 \d p \over
        e^{\beta({p^2\over 2} + V({\bf x})-\mu)} + 1}
  = {\sqrt 2\over\pi^2} \int_0^{\infty} {\sqrt\epsilon \d \epsilon \over
        e^{\beta(\epsilon + V({\bf x})-\mu)} + 1}
  =

  = {\sqrt 2\over\pi^2 \beta^{3\over2}} \int_0^{\infty} {\sqrt y \d y \over
        e^{y - \beta(\mu - V({\bf x}))} + 1}
  = {\sqrt 2\over\pi^2 \beta^{3\over2}}
        I_{1\over2}\left(\beta(\mu - V({\bf x}))\right)

At low temperature (T\to0) we have \beta \to \infty, I_{1\over2}(x) \to {2\over3} x^{3\over 2} and we obtain:

n_e({\bf x}) \to
  {2\sqrt 2\over 3\pi^2 \beta^{3\over2}}
        \left(\beta(\mu - V({\bf x}))\right)^{3\over2}
  ={\left[2(\mu - V({\bf x}))\right]^{3\over2} \over 3\pi^2}

Identical with (8.9.9.1). We can see that the chemical potential \mu becomes the Fermi energy E_f in the limit T\to0. In the finite-temperature case, \mu is determined from the normalization condition for the number of electrons N:

N = \int n_e({\bf x})\, d^3 x

The kinetic energy is

E_{kin} = \int \d^3 x \int 2 {\d^3p\over (2\pi)^3} {p^2\over 2}
    {1\over e^{\beta(E-\mu)}+1}
=

= \int \d^3 x \int_0^\infty 2 {4\pi p^2 \d p\over (2\pi)^3} {p^2\over 2}
    {1\over e^{\beta(E-\mu)}+1} =

= \int \d^3 x \int_0^\infty 2 {4\pi \sqrt 2\sqrt \epsilon \d \epsilon \over
    (2\pi)^3} \epsilon {1\over e^{\beta(\epsilon + V({\bf x})-\mu)}+1} =

= {\sqrt 2 \over \pi^2} \int \d^3 x \int_0^\infty
    {\epsilon^{3\over2} \d \epsilon \over
    e^{\beta(\epsilon + V({\bf x})-\mu)}+1} =

= {\sqrt 2 \over \pi^2 \beta^{5\over2}} \int \d^3 x \int_0^\infty
    {y^{3\over2} \d y \over e^{y - \beta(\mu -V({\bf x}))}+1} =

= {\sqrt 2 \over \pi^2 \beta^{5\over2}}
    \int I_{3\over2}\left(\beta(\mu - V({\bf x}))\right) \d^3 x

From the last formula it can be shown that the kinetic energy is equal to

E_{kin} = {3\over 2} P V - {1\over 2} E_{en} - {1\over 2} E_{ee}

The potential energy is equal to:

E_{pot} = E_{en} + E_{ee}

The internal energy E is equal to:

E = E_{kin} + E_{pot}
    = E_{kin} + E_{en} + E_{ee} =

    = {3\over 2} P V + {1\over 2} E_{en} + {1\over 2} E_{ee}

The entropy S is equal to:

TS
   = -{1\over\beta}
     \sum_i [n_i\log n_i + (1-n_i)\log(1-n_i)] =

   = -{1\over\beta}
     \sum_i \left[n_i\log\left(n_i\over 1-n_i\right)
        + \log(1-n_i)\right] =

   =  \left[\sum_i n_i\epsilon_i\right]
   +
    \left[-\sum_i n_i \mu\right]
   +
        \left[-{1\over\beta} \sum_i\log(1-n_i)\right] =

   = \left[E_{kin} + E_{en} + 2 E_{ee}\right]
    +
    \left[-\mu N\right]
    +
    \left[{2\over3}E_{kin}\right]
    =

   = {5\over3}E_{kin} + E_{en} + 2 E_{ee} -\mu N =

   = {5\over2}P V + {1\over6}E_{en} + {7\over6}E_{ee} -\mu N

where n_i={1\over1+e^{\beta(\epsilon_i-\mu)}} is the number of states at energy \epsilon_i. We used the following calculation expressing one of the sums in terms of the kinetic energy:

-{1\over\beta} \sum_i\log(1-n_i) =

    = -{1\over\beta}\int {2\d^3 x \d^3 p\over (2\pi)^3}
        \log {e^{\beta(E-\mu)}\over 1+e^{\beta(E-\mu)}} =

    = -{\sqrt 2\over \pi^2 \beta^{5\over2}}\int \d^3 x \int_0^\infty
        \sqrt{y}\, \d y
        \log {e^{y-\beta(\mu-V({\bf x}))}\over
            1+e^{y-\beta(\mu-V({\bf x}))}} =

    = -{\sqrt 2\over \pi^2 \beta^{5\over2}}\int \d^3 x \left[
        -{2\over3}\int_0^\infty {y^{3\over2} \d y \over
            1+e^{y-\beta(\mu-V({\bf x}))}} \right] =

    = {2\over 3}{\sqrt 2\over \pi^2 \beta^{5\over2}}\int
        I_{3\over2}\left(\beta(\mu-V({\bf x}))\right) \d^3 x =

    = {2\over 3} E_{kin} \,,

where we used E={p^2\over2}+V({\bf x}).

The free energy is equal to:

F = E - TS = -{2\over3}E_{kin} - E_{ee} + \mu N =

    = -PV + {1\over3}E_{en} - {2\over3}E_{ee} +\mu N

The grand potential is equal to:

\Omega = F - \mu N = -{2\over3}E_{kin} - E_{ee} =

    = -PV + {1\over3}E_{en} - {2\over3}E_{ee}

We can now express the free energy functional F_e[\beta, n_e] as a function of the density:

F_e[\beta, n_e] = -{2\over3}E_{kin} - E_{ee} + \mu N =

    = \int \left(-{2\sqrt2 \over 3 \pi^2 \beta^{5\over2}}
          I_{3\over2}(\Phi(n_e({\bf x})))
          -\half n_e({\bf x}) V_{ee}({\bf x})
        + \mu n_e({\bf x}) \right)\d^3 x =

    = \int \left(-{2\sqrt2 \over 3 \pi^2 \beta^{5\over2}}
          I_{3\over2}(\Phi(n_e({\bf x})))
          -\half n_e({\bf x}) V_{ee}({\bf x})
        + {1\over \beta} n_e({\bf x}) \Phi(n_e({\bf x}))
            + n_e({\bf x}) V({\bf x}) \right)\d^3 x =

    = \int \left(-{2\sqrt2 \over 3 \pi^2 \beta^{5\over2}}
          I_{3\over2}(\Phi(n_e({\bf x})))
          -\half n_e({\bf x}) V_{ee}({\bf x})
        + {1\over \beta} n_e({\bf x}) \Phi(n_e({\bf x}))
            + n_e({\bf x}) (V_{en}({\bf x}) + V_{ee}({\bf x})
                + V_{xc}({\bf x}))
              \right)\d^3 x =

    = \int \left(-{2\sqrt2 \over 3 \pi^2 \beta^{5\over2}}
          I_{3\over2}(\Phi(n_e({\bf x})))
        + {1\over \beta} n_e({\bf x}) \Phi(n_e({\bf x}))
            + n_e({\bf x}) (V_{en}({\bf x}) +\half V_{ee}({\bf x})
                + V_{xc}({\bf x}))
              \right)\d^3 x =

    = \left( -{2\over3}E_{kin}
        + \int {1\over \beta} n_e({\bf x}) \Phi(n_e({\bf x}))\, \d^3 x
          \right)
        + E_{en} + E_{ee} + E_{xc}

8.9.10. Orbital Free Density Functional Theory

The orbital-free electronic free energy is given by:

F_e[n_e] = T_0[n_e] + U_{en}[n_e] + U_{ee}[n_e] + F_{xc}[n_e]\,,

where the kinetic energy can be written in a few different equivalent ways as

T_0[n_e] =

= {\sqrt 2\over\pi^2 \beta^{5\over2}}
  \int \left(
            I_{1\over2}\left(\beta(\mu - V({\bf x}))\right)
\beta(\mu-V({\bf x})) -
{2\over 3} I_{3\over2}(\beta(\mu-V({\bf x})))
\right) \d^3 x =

= {1\over\beta} \int \left(
n_e(\mathbf{x}) \beta(\mu-V({\bf x})) -
{2\sqrt2\over 3\pi^2 \beta^{3\over2}} I_{3\over2}(\beta(\mu-V({\bf x})))
\right) \d^3 x =

= {1\over\beta} \int \left(
n_e(\mathbf{x}) \Phi(n_e(\mathbf{x})) -
{2\sqrt2\over 3\pi^2 \beta^{3\over2}} I_{3\over2}(\Phi(n_e(\mathbf{x})))
\right) \d^3 x =

= {1\over\beta} \int \left(
n_e(\mathbf{x}) I_{1\over2}^{-1}(y) -
{2n_e(\mathbf{x})\over 3 y} I_{3\over2}(I_{1\over2}^{-1}(y))
\right) \d^3 x =

= {1\over\beta} \int
n_e(\mathbf{x}) \left( I_{1\over2}^{-1}(y) -
{2\over 3 y} I_{3\over2}(I_{1\over2}^{-1}(y))\right)
\d^3 x =

={1\over\beta} \int n_e(\mathbf{x}) f(y) \d^3 x\,;\quad\quad
y={\pi^2\over\sqrt2}\beta^{3\over2} n_e\,,

where f(y) is a special function of one variable, composed of a Fermi-Dirac Integral of order 3\over2 and its inverse of order 1\over2:

f(y) = I_{1\over2}^{-1}(y) - {2\over 3 y} I_{3\over2}(I_{1\over2}^{-1}(y))

the electron-nuclei term has the form

U_{en}[n_e]
    = \int {n_e(\mathbf{x}) n_n(\mathbf{x}')\over
        |\mathbf{x} - \mathbf{x}'|}\d^3 x\d^3 x'
    = \int n_e(\mathbf{x}) V_{en}(\mathbf{x}) \d^3 x\,,

The electron-electron (Hartree) term takes the form:

U_{ee}[n_e]
    = \half \int {n_e(\mathbf{x}) n_e(\mathbf{x}')\over
        |\mathbf{x} - \mathbf{x}'|}\d^3 x\d^3 x'
    = \half \int n_e(\mathbf{x}) V_{ee}(\mathbf{x}) \d^3 x\,,

and the exchange and correlation functional F_{xc}[n_e] is given by the Perdew-Zunger LDA:

F_{xc}[n_e]
    = \int n_e(\mathbf{x}) \epsilon_{xc}^{LD}(n_e) \d^3 x\,,

n_e(\mathbf{x}) is the (positive) electron density, n_n(\mathbf{x}) is the (positive) nuclei density.

We minimize this free energy under the condition of particle conservation. The constrained functional is (we use n\equiv n_e from now on):

\Omega[n] = F_e[n] - \epsilon \left(\int n(\mathbf{x}) \d^3 x - N\right)

The variational solution is:

{\delta \Omega[n] \over \delta n} = 0

Or:

(8.9.10.1){\delta F_e[n] \over \delta n} = \epsilon

Finally we get:

(8.9.10.2)H[n] \equiv {\delta F_e[n] \over \delta n} =
{\delta T_0[n] \over \delta n}
+ {\delta U_{en}[n] \over \delta n}
+ {\delta U_{ee}[n] \over \delta n}
+ {\delta F_{xc}[n] \over \delta n}
= \epsilon

The individual terms are:

{\delta T_0[n] \over \delta n} =
    {1\over\beta} \left( f(y) + n(\mathbf{x}) {\d f(y) \over \d y}
    {\pi^2\over\sqrt2}\beta^{3\over2} \right)

and

{\delta U_{en}[n] \over \delta n}
    = \int {n_n(\mathbf{x}')\over
        |\mathbf{x} - \mathbf{x}'|}\d^3 x'
    \equiv V_{en}(\mathbf{x})

and

{\delta U_{ee}[n] \over \delta n}
    = \int {n(\mathbf{x}')\over
        |\mathbf{x} - \mathbf{x}'|}\d^3 x'
    \equiv V_{ee}(\mathbf{x})

and

{\delta F_{xc}[n] \over \delta n}
    = \epsilon_{xc}^{LD}(n) + n(\mathbf{x})
        {\d \epsilon_{xc}^{LD}(n) \over \d n}
    \equiv V_{xc}(\mathbf{x})

All together the Hamiltonian is:

H[n] =
    {1\over\beta} \left( f(y) + n(\mathbf{x}) {\d f(y) \over \d y}
    {\pi^2\over\sqrt2}\beta^{3\over2} \right)
    +
    V_{en}(\mathbf{x})
    +
    V_{ee}(\mathbf{x})
    +
    V_{xc}(\mathbf{x})

We can also introduce an artificial orbital \psi(\mathbf{x}) as follows:

n(\mathbf{x}) = \psi^2(\mathbf{x})

and minimize \Omega with respect to \psi:

{\delta \Omega[n] \over \delta \psi} = 0 \,.

We will use tilde to denote functions in terms of \psi. So \Omega[n] =
\Omega[\psi^2] = \tilde \Omega[\psi]. Using the relation

{\delta \over \delta \psi}
    = {\d n \over \d \psi} {\delta \over \delta n}
    = 2 \psi {\delta \over \delta n}

we obtain:

{\delta \Omega[n] \over \delta \psi}
    = 2 \psi {\delta \Omega[n] \over \delta n} = 0 \,.

So the equation (8.9.10.1) gets multiplied by 2\psi:

2\psi {\delta F_e[n] \over \delta n} = 2\epsilon \psi

as well as the equation (8.9.10.2):

\tilde H[\psi] \equiv {\delta \tilde F_e[\psi] \over \delta \psi}
    = {\delta F_e[n] \over \delta \psi}
    = 2 \psi {\delta F_e[n] \over \delta n}
    = 2 H[n] \psi
    = 2\epsilon \psi

So the Hamiltonian H[n] expressed using n and the Hamiltonian \tilde H[\psi] expressed using \psi are related by \tilde H[\psi] = 2 H[n] \psi.

Free Energy Minimization

For clarity, we will be using H[n] from equation (8.9.10.2) as our main quantity, but we will also write the final relations using \tilde H[\psi] for completeness.

We start with some initial guess for \ket{\psi} (it must be normalized as \braket{\psi | \psi}=N). Let’s calculate \epsilon:

H[n] = \epsilon

H[n]\ket{\psi} = \epsilon\ket{\psi}

\braket{\psi | H[n] | \psi} = \epsilon\braket{\psi | \psi}

\braket{\psi | H[n] | \psi} = \epsilon N

\epsilon = {1\over N} \braket{\psi | H[n] | \psi}
    \equiv {1\over N} \int H[n] \psi^2(\mathbf{x}) \d^3 x
    = {1\over 2N} \int \tilde H[\psi] \psi(\mathbf{x}) \d^3 x

We calculate the steepest-descent (SD) vector \ket{\chi}:

\ket{\chi} = 2(\epsilon - H[n])\ket{\psi}
    \equiv 2\epsilon \ket{\psi} - \ket{\tilde H[\psi]}

The conjugate-gradient (CG) vector \ket{\varphi} is calculated as:

\ket{\varphi} = \ket{\chi} + {\braket{\chi | \chi} \over
    \braket{\chi_{k-1} | \chi_{k-1}}} \ket{\varphi}

To satisfy the normalization constraint of \ket{\psi}, the CG vector is further orthogonalized to \ket{\psi} and normalized to N (this step is one particular, but not the only way to impose the normalization constraint):

\ket{\varphi'} = \left(1 - {1\over N} \ket{\psi}\bra{\psi}\right)
    \ket{\varphi}

\ket{\varphi''} = \sqrt{N \over \braket{\varphi' | \varphi'}} \,
    \ket{\varphi'}

That is, now \braket{\varphi'' | \psi}=0 and \braket{\varphi'' | \varphi''} = N. The new CG vector \ket{\psi_{k+1}} is then updated as usual in CG by \ket{\psi} + \alpha \ket{\varphi''}, but then it must be normalized. As such, equivalently, it is updated by a linear combination of \ket{\psi} and \ket{\varphi''}:

\ket{\psi_{k+1}} = a \ket{\psi} + b \ket{\varphi''}

such that it remains normalized:

\braket{\psi_{k+1} | \psi_{k+1}}
    = (a \bra{\psi} + b \bra{\varphi''})
    (a \ket{\psi} + b \ket{\varphi''})
    = (a^2 + b^2) N = N

So a, b are any real numbers satisfying the equation a^2+b^2=1, whose parametric solution is a=\cos\theta, b=\sin\theta with 0 \le \theta <
2\pi:

\ket{\psi_{k+1}} = \cos\theta \ket{\psi} + \sin\theta \ket{\varphi''}

where \theta is determined by minimizing the free energy F_e[\psi_{k+1}] as a function of \theta.

8.9.11. References

[DFT](1, 2, 3)
    1. Dreizler, E. K. U. Gross: Density functional theory: an approach to the quantum many-body problem
[pickett](1, 2)
    1. Pickett, Pseudopotential methods in condensed matter applications, Computer Physics reports, Volume 9, Issue 3, April 1989, Pages 115-197, ISSN 0167-7977, DOI: 10.1016/0167-7977(89)90002-6. (http://www.sciencedirect.com/science/article/B6X3V-46R02CR-1J/2/804d9ecaa49469aa5e1050dc007f4a61)