8.2: Boltzmann Transport Theory
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Derivation of the Boltzmann equation
For simplicity of presentation, we assume point particles. Recall that f(r,p,t)d3rd3p≡{\rm\# of particles with positions within d3r ofr and momenta within d3p of p at time t. We now ask how the distribution functions f(r,p,t) evolves in time. It is clear that in the absence of collisions, the distribution function must satisfy the continuity equation, ∂f∂t+∇⋅(uf)=0 . This is just the condition of number conservation for particles. Take care to note that ∇ and u are six-dimensional phase space vectors: u=( ˙x , ˙y , ˙z , ˙px , ˙py , ˙pz )∇=(∂∂x,∂∂y,∂∂z,∂∂px,∂∂py,∂∂pz) . The continuity equation describes a distribution in which each constituent particle evolves according to a prescribed dynamics, which for a mechanical system is specified by drdt=∂H∂p=v(p),dpdt=−∂H∂r=F∗ext , where F is an external applied force. Here, H(p,r)=ε(p)+U∗ext(r) . For example, if the particles are under the influence of gravity, then Uext(r)=mg⋅r and F=−∇U∗ext=−mg.
Note that as a consequence of the dynamics, we have ∇⋅u=0, phase space flow is incompressible, provided that ε(p) is a function of p alone, and not of r. Thus, in the absence of collisions, we have ∂f∂t+u⋅∇f=0 . The differential operator Dt≡∂t+u⋅∇ is sometimes called the ‘convective derivative’, because D∗tf is the time derivative of f in a comoving frame of reference.
Next we must consider the effect of collisions, which are not accounted for by the semiclassical dynamics. In a collision process, a particle with momentum p and one with momentum ˜p can instantaneously convert into a pair with momenta p′ and ˜p′, provided total momentum is conserved: p+˜p=p′+˜p′. This means that Dtf≠0. Rather, we should write ∂f∂t+˙r⋅∂f∂r+˙p⋅∂f∂p=(∂f∂t)†coll where the right side is known as the collision integral. The collision integral is in general a function of r, p, and t and a functional of the distribution f.
After a trivial rearrangement of terms, we can write the Boltzmann equation as ∂f∂t=(∂f∂t)†str+(∂f∂t)†coll , where (∂f∂t)†str≡−˙r⋅∂f∂r−˙p⋅∂f∂p is known as the streaming term. Thus, there are two contributions to ∂f/∂t : streaming and collisions.
Collisionless Boltzmann equation
In the absence of collisions, the Boltzmann equation is given by ∂f∂t+∂ε∂p⋅∂f∂r−∇U∗ext⋅∂f∂p=0 . In order to gain some intuition about how the streaming term affects the evolution of the distribution f(r,p,t), consider a case where F∗ext=0. We then have ∂f∂t+pm⋅∂f∂r=0 . Clearly, then, any function of the form f(r,p,t)=φ(r−v(p)t,p) will be a solution to the collisionless Boltzmann equation, where v(p)=∂ε∂p. One possible solution would be the Boltzmann distribution, f(r,p,t)=eμ/kBTe−p2/2mkBT , which is time-independent1. Here we have assumed a ballistic dispersion, ε(p)=p2/2m.
For a slightly less trivial example, let the initial distribution be φ(r,p)=Ae−r2/2σ2e−p2/2κ2, so that f(r,p,t)=Ae−(r−ptm)2/2σ2e−p2/2κ2 . Consider the one-dimensional version, and rescale position, momentum, and time so that f(x,p,t)=Ae−12(ˉx−ˉpˉt)2e−12ˉp2 . Consider the level sets of f, where f(x,p,t)=Ae−12α2. The equation for these sets is ˉx=ˉpˉt±√α2−ˉp2 . For fixed ˉt, these level sets describe the loci in phase space of equal probability densities, with the probability density decreasing exponentially in the parameter α2. For ˉt=0, the initial distribution describes a Gaussian cloud of particles with a Gaussian momentum distribution. As ˉt increases, the distribution widens in ˉx but not in ˉp – each particle moves with a constant momentum, so the set of momentum values never changes. However, the level sets in the (ˉx,ˉp) plane become elliptical, with a semimajor axis oriented at an angle θ=ctn−1(t) with respect to the ˉx axis. For ˉt>0, he particles at the outer edges of the cloud are more likely to be moving away from the center. See the sketches in Figure [Fstreaming]
Suppose we add in a constant external force F∗ext. Then it is easy to show (and left as an exercise to the reader to prove) that any function of the form f(r,p,t)=Aφ(r−ptm+F∗extt22m,p−F∗exttm) satisfies the collisionless Boltzmann equation (ballistic dispersion assumed).
Collisional invariants
Consider a function A(r,p) of position and momentum. Its average value at time t is A(t)=∫d3rd3pA(r,p)f(r,p,t) . Taking the time derivative, dAdt=∫d3rd3pA(r,p)∂f∂t=∫d3rd3pA(r,p){−∂∂r⋅(˙rf)−∂∂p⋅(˙pf)+(∂f∂t)†coll}=∫d3rd3p{(∂A∂r⋅drdt+∂A∂p⋅dpdt)f+A(r,p)(∂f∂t)†coll} . Hence, if A is preserved by the dynamics between collisions, then2 dAdt=∂A∂r⋅drdt+∂A∂p⋅dpdt=0 . We therefore have that the rate of change of A is determined wholly by the collision integral dAdt=∫d3rd3pA(r,p)(∂f∂t)†coll . Quantities which are then conserved in the collisions satisfy ˙A=0. Such quantities are called collisional invariants. Examples of collisional invariants include the particle number (A=1), the components of the total momentum (A=p∗μ) (in the absence of broken translational invariance, due to the presence of walls), and the total energy (A=ε(p)).
Scattering processes
What sort of processes contribute to the collision integral? There are two broad classes to consider. The first involves potential scattering, where a particle in state |Γ⟩ scatters, in the presence of an external potential, to a state |Γ′⟩. Recall that Γ is an abbreviation for the set of kinematic variables, Γ=(p,L) in the case of a diatomic molecule. For point particles, Γ=(p∗x,p∗y,p∗z) and dΓ=d3p.
We now define the function w(Γ′|Γ) such that w(Γ′|Γ)f(r,Γ;t)dΓdΓ′={rate at which a particle within dΓ of (r,Γ)scatters to within dΓ′ of (r,Γ′) at time t. The units of wdΓ are therefore 1/T. The differential scattering cross section for particle scattering is then dσ=w(Γ′|Γ)n|v|dΓ′ , where v=p/m is the particle’s velocity and n the density.
The second class is that of two-particle scattering processes, |ΓΓ∗1⟩→|Γ′Γ′1⟩. We define the scattering function w(Γ′Γ′1|ΓΓ∗1) by w(Γ′Γ′1|ΓΓ∗1)f∗2(r,Γ;r,Γ∗1;t)dΓdΓ∗1dΓ′dΓ′1={rate at which two particles within dΓ of (r,Γ)and within dΓ∗1 of (r,Γ∗1) scatter into states withindΓ′ of (r,Γ′) and dΓ′1 of (r,Γ′1) at time t\,, where f∗2(r,p;r′,p′;t)=⟨∑i,jδ(x∗i(t)−r)δ(p∗i(t)−p)δ(x∗j(t)−r′)δ(p∗j(t)−p′)⟩ is the nonequilibrium two-particle distribution for point particles. The differential scattering cross section is dσ=w(Γ′Γ′1|ΓΓ∗1)|v−v∗1|dΓ′dΓ′1 .
We assume, in both cases, that any scattering occurs locally, the particles attain their asymptotic kinematic states on distance scales small compared to the mean interparticle separation. In this case we can treat each scattering process independently. This assumption is particular to rarefied systems, gases, and is not appropriate for dense liquids. The two types of scattering processes are depicted in Figure [FCIscatt].
In computing the collision integral for the state |r,Γ⟩, we must take care to sum over contributions from transitions out of this state, |Γ⟩→|Γ′⟩, which reduce f(r,Γ), and transitions into this state, |Γ′⟩→|Γ⟩, which increase f(r,Γ). Thus, for one-body scattering, we have DDtf(r,Γ;t)=(∂f∂t)†coll=∫dΓ′{w(Γ|Γ′)f(r,Γ′;t)−w(Γ′|Γ)f(r,Γ;t)} . For two-body scattering, we have DDtf(r,Γ;t)=(∂f∂t)†coll=∫dΓ∗1∫dΓ′∫dΓ′1{w(ΓΓ∗1|Γ′Γ′1)f∗2(r,Γ′;r,Γ′1;t)−w(Γ′Γ′1|ΓΓ∗1)f∗2(r,Γ;r,Γ∗1;t)} . Unlike the one-body scattering case, the kinetic equation for two-body scattering does not close, since the LHS involves the one-body distribution f≡f∗1 and the RHS involves the two-body distribution f∗2. To close the equations, we make the approximation f∗2(r,Γ′;˜r,˜Γ;t)≈f(r,Γ;t)f(˜r,˜Γ;t) . We then have DDtf(r,Γ;t)=∫dΓ∗1∫dΓ′∫dΓ′1{w(ΓΓ∗1|Γ′Γ′1)f(r,Γ′;t)f(r,Γ′1;t)−w(Γ′Γ′1|ΓΓ∗1)f(r,Γ;t)f(r,Γ∗1;t)} .
Detailed balance
Classical mechanics places some restrictions on the form of the kernel w(ΓΓ∗1|Γ′Γ′1). In particular, if \boldsymbol{\Gamma^\sss{T}=(-\Bp,-\BL)} denotes the kinematic variables under time reversal, then \boldsymbol{w\big(\Gamma'\Gamma'_1 \, | \, \Gamma\Gamma_1\big)= w\big(\Gamma^\sss{T}\Gamma^\sss{T}_1 \, | \,\Gamma'{}^\sss{T}\Gamma'_1{}^\sss{T}\big)\ . \label{TRw}} This is because the time reverse of the process |ΓΓ∗1⟩→|Γ′Γ′1⟩ is \boldsymbol{\tket{\Gamma'{}^\ssr{T}\Gamma'_1{}^\ssr{T}}\to\tket{\Gamma^\ssr{T}\Gamma^\ssr{T}_1}}.
In equilibrium, we must have \boldsymbol{w\big(\Gamma'\Gamma'_1 \, | \, \Gamma\Gamma_1\big)\,\,f^0(\Gamma)\,f^0(\Gamma\ns_1) \,d^4\!\Gamma= w\big(\Gamma^\sss{T}\Gamma^\sss{T}_1 \, | \,\Gamma'{}^\sss{T}\Gamma'_1{}^\sss{T}\big)\,f^0(\Gamma'{}^\sss{T})\,f^0(\Gamma'_1{}^\sss{T}) \,d^4\!\Gamma^\sss{T}} where \boldsymbol{d^4\!\Gamma\equiv d\Gamma\,d\Gamma\ns_1\,d\Gamma' d\Gamma'_1\qquad,\qquad d^4\!\Gamma^\sss{T}\equiv d\Gamma^\sss{T}\,d\Gamma_1^\sss{T}\,d\Gamma'{}^\sss{T} d\Gamma'_1{}^\sss{T}\ .} Since \boldsymbol{d\Gamma=d\Gamma^\sss{T}} , we may cancel the differentials above, and after invoking Equation [TRw] and suppressing the common r label, we find \boldsymbol{f^0(\Gamma)\,f^0(\Gamma\ns_1)=f^0(\Gamma'{}^\sss{T})\,f^0(\Gamma'_1{}^\sss{T})\ .} This is the condition of detailed balance. For the Boltzmann distribution, we have f0(Γ)=Ae−ε/kBT , where A is a constant and where ε=ε(Γ) is the kinetic energy, ε(Γ)=p2/2m in the case of point particles. Note that \boldsymbol{\ve({\Gamma^\sss{T}})=\ve(\Gamma)}. Detailed balance is satisfied because the kinematics of the collision requires energy conservation: ε+ε∗1=ε′+ε′1 . Since momentum is also kinematically conserved, p+p∗1=p′+p′1 , any distribution of the form f0(Γ)=Ae−(ε−p⋅V)/kBT also satisfies detailed balance, for any velocity parameter V. This distribution is appropriate for gases which are flowing with average particle V.
In addition to time-reversal, parity is also a symmetry of the microscopic mechanical laws. Under the parity operation P, we have r→−r and p→−p. Note that a pseudovector such as L=r×p is unchanged under P. Thus, \boldsymbol{\Gamma^\sss{P}=(-\Bp,\BL)}. Under the combined operation of C=PT, we have \boldsymbol{\Gamma^\sss{C}=(\Bp,-\BL)}. If the microscopic Hamiltonian is invariant under C, then we must have \boldsymbol{w\big(\Gamma'\Gamma'_1 \, | \, \Gamma\Gamma\ns_1\big)=w\big(\Gamma^\sss{C}\Gamma_1^\sss{C} \, | \, \Gamma'{}^\sss{C}\Gamma'_1{}^\sss{C}\big)\ .} For point particles, invariance under T and P then means w(p′,p′1|p,p∗1)=w(p,p∗1|p′,p′1) , and therefore the collision integral takes the simplified form, Df(p)Dt=(∂f∂t)†coll=∫d3p∗1∫d3p′∫d3p′1w(p′,p′1|p,p∗1){f(p′)f(p′1)−f(p)f(p∗1)} , where we have suppressed both r and t variables.
The most general statement of detailed balance is f0(Γ′)f0(Γ′1)f0(Γ)f0(Γ∗1)=w(Γ′Γ′1|ΓΓ∗1)w(ΓΓ∗1|Γ′Γ′1) . Under this condition, the collision term vanishes for f=f0, which is the equilibrium distribution.
Kinematics and cross section
We can rewrite Equation [BEwp] in the form Df(p)Dt=∫d3p∗1∫dΩ|v−v∗1|∂σ∂Ω{f(p′)f(p′1)−f(p)f(p∗1)} , where ∂σ∂Ω is the differential scattering cross section. If we recast the scattering problem in terms of center-of-mass and relative coordinates, we conclude that the total momentum is conserved by the collision, and furthermore that the energy in the CM frame is conserved, which means that the magnitude of the relative momentum is conserved. Thus, we may write p′−p′1=|p−p∗1|ˆΩ, where ˆΩ is a unit vector. Then p′ and p′1 are determined to be p′=12(p+p∗1+|p−p∗1|ˆΩ)p′1=12(p+p∗1−|p−p∗1|ˆΩ) .
H-theorem
Let’s consider the Boltzmann equation with two particle collisions. We define the local ( r-dependent) quantity ρ∗φ(r,t)≡∫dΓφ(Γ,f)f(Γ,r,t) . At this point, φ(Γ,f) is arbitrary. Note that the φ(Γ,f) factor has r and t dependence through its dependence on f, which itself is a function of r, Γ, and t. We now compute ∂ρ∗φ∂t=∫dΓ∂(φf)∂t=∫dΓ∂(φf)∂f∂f∂t=−∫dΓu⋅∇(φf)−∫dΓ∂(φf)∂f(∂f∂t)†coll=−∮dΣˆn⋅(uφf)−∫dΓ∂(φf)∂f(∂f∂t)†coll . The first term on the last line follows from the divergence theorem, and vanishes if we assume f=0 for infinite values of the kinematic variables, which is the only physical possibility. Thus, the rate of change of ρ∗φ is entirely due to the collision term. Thus, ∂ρ∗φ∂t=∫dΓ∫dΓ∗1∫dΓ′∫dΓ′1{w(Γ′Γ′1|ΓΓ∗1)ff∗1χ−w(ΓΓ∗1|Γ′Γ′1)f′f′1χ}=∫dΓ∫dΓ∗1∫dΓ′∫dΓ′1w(Γ′Γ′1|ΓΓ∗1)ff∗1(χ−χ′) , where f≡f(Γ), f′≡f(Γ′), f∗1≡f(Γ∗1), f′1≡f(Γ′1), χ=χ(Γ), with χ=∂(φf)∂f=φ+f∂φ∂f . We now invoke the symmetry w(Γ′Γ′1|ΓΓ∗1)=w(Γ′1Γ′|Γ∗1Γ) , which allows us to write ∂ρ∗φ∂t=12∫dΓ∫dΓ∗1∫dΓ′∫dΓ′1w(Γ′Γ′1|ΓΓ∗1)ff∗1(χ+χ∗1−χ′−χ′1) . This shows that ρ∗φ is preserved by the collision term if χ(Γ) is a collisional invariant.
Now let us consider φ(f)=lnf. We define h≡ρ|φ=lnf. We then have ∂h∂t=−12∫dΓ∫dΓ∗1∫dΓ′∫dΓ′1wf′f′1⋅xlnx , where w≡w(Γ′Γ′1|ΓΓ∗1) and x≡ff∗1/f′f′1. We next invoke the result ∫dΓ′∫dΓ′1w(Γ′Γ′1|ΓΓ∗1)=∫dΓ′∫dΓ′1w(ΓΓ∗1|Γ′Γ′1) which is a statement of unitarity of the scattering matrix3. Multiplying both sides by f(Γ)f(Γ∗1), then integrating over Γ and Γ∗1, and finally changing variables (Γ,Γ∗1)↔(Γ′,Γ′1), we find 0=∫dΓ∫dΓ∗1∫dΓ′∫dΓ′1w(ff∗1−f′f′1)=∫dΓ∫dΓ∗1∫dΓ′∫dΓ′1wf′f′1(x−1) . Multiplying this result by 12 and adding it to the previous equation for ˙h, we arrive at our final result, ∂h∂t=−12∫dΓ∫dΓ∗1∫dΓ′∫dΓ′1wf′f′1(xlnx−x+1) . Note that w, f′, and f′1 are all nonnegative. It is then easy to prove that the function g(x)=xlnx−x+1 is nonnegative for all positive x values4, which therefore entails the important result ∂h(r,t)∂t≤0 . Boltzmann’s H function is the space integral of the h density: H=∫d3rh.
Thus, everywhere in space, the function h(r,t) is monotonically decreasing or constant, due to collisions. In equilibrium, ˙h=0 everywhere, which requires x=1, f0(Γ)f0(Γ∗1)=f0(Γ′)f0(Γ′1) , or, taking the logarithm, lnf0(Γ)+lnf0(Γ∗1)=lnf0(Γ′)+lnf0(Γ′1) . But this means that lnf0 is itself a collisional invariant, and if 1, p, and ε are the only collisional invariants, then lnf0 must be expressible in terms of them. Thus, lnf0=μkBT+V⋅pkBT−εkBT , where μ, V, and T are constants which parameterize the equilibrium distribution f0(p), corresponding to the chemical potential, flow velocity, and temperature, respectively.