31 |
|
$F_{ji}$ be the force that particle $j$ exerts on particle $i$. |
32 |
|
Newton's third law states that |
33 |
|
\begin{equation} |
34 |
< |
F_{ij} = -F_{ji} |
34 |
> |
F_{ij} = -F_{ji}. |
35 |
|
\label{introEquation:newtonThirdLaw} |
36 |
|
\end{equation} |
37 |
– |
|
37 |
|
Conservation laws of Newtonian Mechanics play very important roles |
38 |
|
in solving mechanics problems. The linear momentum of a particle is |
39 |
|
conserved if it is free or it experiences no force. The second |
62 |
|
\end{equation} |
63 |
|
If there are no external torques acting on a body, the angular |
64 |
|
momentum of it is conserved. The last conservation theorem state |
65 |
< |
that if all forces are conservative, Energy |
66 |
< |
\begin{equation}E = T + V \label{introEquation:energyConservation} |
65 |
> |
that if all forces are conservative, energy is conserved, |
66 |
> |
\begin{equation}E = T + V. \label{introEquation:energyConservation} |
67 |
|
\end{equation} |
68 |
< |
is conserved. All of these conserved quantities are |
69 |
< |
important factors to determine the quality of numerical integration |
70 |
< |
schemes for rigid bodies \cite{Dullweber1997}. |
68 |
> |
All of these conserved quantities are important factors to determine |
69 |
> |
the quality of numerical integration schemes for rigid bodies |
70 |
> |
\cite{Dullweber1997}. |
71 |
|
|
72 |
|
\subsection{\label{introSection:lagrangian}Lagrangian Mechanics} |
73 |
|
|
74 |
|
Newtonian Mechanics suffers from two important limitations: motions |
75 |
< |
can only be described in cartesian coordinate systems. Moreover, It |
76 |
< |
become impossible to predict analytically the properties of the |
75 |
> |
can only be described in cartesian coordinate systems. Moreover, it |
76 |
> |
becomes impossible to predict analytically the properties of the |
77 |
|
system even if we know all of the details of the interaction. In |
78 |
|
order to overcome some of the practical difficulties which arise in |
79 |
|
attempts to apply Newton's equation to complex system, approximate |
84 |
|
|
85 |
|
Hamilton introduced the dynamical principle upon which it is |
86 |
|
possible to base all of mechanics and most of classical physics. |
87 |
< |
Hamilton's Principle may be stated as follows, |
88 |
< |
|
89 |
< |
The actual trajectory, along which a dynamical system may move from |
90 |
< |
one point to another within a specified time, is derived by finding |
91 |
< |
the path which minimizes the time integral of the difference between |
93 |
< |
the kinetic, $K$, and potential energies, $U$. |
87 |
> |
Hamilton's Principle may be stated as follows: the actual |
88 |
> |
trajectory, along which a dynamical system may move from one point |
89 |
> |
to another within a specified time, is derived by finding the path |
90 |
> |
which minimizes the time integral of the difference between the |
91 |
> |
kinetic, $K$, and potential energies, $U$, |
92 |
|
\begin{equation} |
93 |
< |
\delta \int_{t_1 }^{t_2 } {(K - U)dt = 0} , |
93 |
> |
\delta \int_{t_1 }^{t_2 } {(K - U)dt = 0}. |
94 |
|
\label{introEquation:halmitonianPrinciple1} |
95 |
|
\end{equation} |
98 |
– |
|
96 |
|
For simple mechanical systems, where the forces acting on the |
97 |
|
different parts are derivable from a potential, the Lagrangian |
98 |
|
function $L$ can be defined as the difference between the kinetic |
135 |
|
p_i = \frac{{\partial L}}{{\partial q_i }} |
136 |
|
\label{introEquation:generalizedMomentaDot} |
137 |
|
\end{equation} |
141 |
– |
|
138 |
|
With the help of the generalized momenta, we may now define a new |
139 |
|
quantity $H$ by the equation |
140 |
|
\begin{equation} |
142 |
|
\label{introEquation:hamiltonianDefByLagrangian} |
143 |
|
\end{equation} |
144 |
|
where $ \dot q_1 \ldots \dot q_f $ are generalized velocities and |
145 |
< |
$L$ is the Lagrangian function for the system. |
146 |
< |
|
151 |
< |
Differentiating Eq.~\ref{introEquation:hamiltonianDefByLagrangian}, |
152 |
< |
one can obtain |
145 |
> |
$L$ is the Lagrangian function for the system. Differentiating |
146 |
> |
Eq.~\ref{introEquation:hamiltonianDefByLagrangian}, one can obtain |
147 |
|
\begin{equation} |
148 |
|
dH = \sum\limits_k {\left( {p_k d\dot q_k + \dot q_k dp_k - |
149 |
|
\frac{{\partial L}}{{\partial q_k }}dq_k - \frac{{\partial |
150 |
|
L}}{{\partial \dot q_k }}d\dot q_k } \right)} - \frac{{\partial |
151 |
|
L}}{{\partial t}}dt \label{introEquation:diffHamiltonian1} |
152 |
|
\end{equation} |
153 |
< |
Making use of Eq.~\ref{introEquation:generalizedMomenta}, the |
154 |
< |
second and fourth terms in the parentheses cancel. Therefore, |
153 |
> |
Making use of Eq.~\ref{introEquation:generalizedMomenta}, the second |
154 |
> |
and fourth terms in the parentheses cancel. Therefore, |
155 |
|
Eq.~\ref{introEquation:diffHamiltonian1} can be rewritten as |
156 |
|
\begin{equation} |
157 |
|
dH = \sum\limits_k {\left( {\dot q_k dp_k - \dot p_k dq_k } |
174 |
|
t}} |
175 |
|
\label{introEquation:motionHamiltonianTime} |
176 |
|
\end{equation} |
177 |
< |
|
184 |
< |
Eq.~\ref{introEquation:motionHamiltonianCoordinate} and |
177 |
> |
where Eq.~\ref{introEquation:motionHamiltonianCoordinate} and |
178 |
|
Eq.~\ref{introEquation:motionHamiltonianMomentum} are Hamilton's |
179 |
|
equation of motion. Due to their symmetrical formula, they are also |
180 |
|
known as the canonical equations of motions \cite{Goldstein2001}. |
188 |
|
statistical mechanics and quantum mechanics, since it treats the |
189 |
|
coordinate and its time derivative as independent variables and it |
190 |
|
only works with 1st-order differential equations\cite{Marion1990}. |
198 |
– |
|
191 |
|
In Newtonian Mechanics, a system described by conservative forces |
192 |
< |
conserves the total energy \ref{introEquation:energyConservation}. |
193 |
< |
It follows that Hamilton's equations of motion conserve the total |
194 |
< |
Hamiltonian. |
192 |
> |
conserves the total energy |
193 |
> |
(Eq.~\ref{introEquation:energyConservation}). It follows that |
194 |
> |
Hamilton's equations of motion conserve the total Hamiltonian. |
195 |
|
\begin{equation} |
196 |
|
\frac{{dH}}{{dt}} = \sum\limits_i {\left( {\frac{{\partial |
197 |
|
H}}{{\partial q_i }}\dot q_i + \frac{{\partial H}}{{\partial p_i |
219 |
|
momentum variables. Consider a dynamic system of $f$ particles in a |
220 |
|
cartesian space, where each of the $6f$ coordinates and momenta is |
221 |
|
assigned to one of $6f$ mutually orthogonal axes, the phase space of |
222 |
< |
this system is a $6f$ dimensional space. A point, $x = (q_1 , \ldots |
223 |
< |
,q_f ,p_1 , \ldots ,p_f )$, with a unique set of values of $6f$ |
224 |
< |
coordinates and momenta is a phase space vector. |
225 |
< |
|
222 |
> |
this system is a $6f$ dimensional space. A point, $x = (\rightarrow |
223 |
> |
q_1 , \ldots ,\rightarrow q_f ,\rightarrow p_1 , \ldots ,\rightarrow |
224 |
> |
p_f )$, with a unique set of values of $6f$ coordinates and momenta |
225 |
> |
is a phase space vector. |
226 |
|
%%%fix me |
227 |
< |
A microscopic state or microstate of a classical system is |
228 |
< |
specification of the complete phase space vector of a system at any |
237 |
< |
instant in time. An ensemble is defined as a collection of systems |
238 |
< |
sharing one or more macroscopic characteristics but each being in a |
239 |
< |
unique microstate. The complete ensemble is specified by giving all |
240 |
< |
systems or microstates consistent with the common macroscopic |
241 |
< |
characteristics of the ensemble. Although the state of each |
242 |
< |
individual system in the ensemble could be precisely described at |
243 |
< |
any instance in time by a suitable phase space vector, when using |
244 |
< |
ensembles for statistical purposes, there is no need to maintain |
245 |
< |
distinctions between individual systems, since the numbers of |
246 |
< |
systems at any time in the different states which correspond to |
247 |
< |
different regions of the phase space are more interesting. Moreover, |
248 |
< |
in the point of view of statistical mechanics, one would prefer to |
249 |
< |
use ensembles containing a large enough population of separate |
250 |
< |
members so that the numbers of systems in such different states can |
251 |
< |
be regarded as changing continuously as we traverse different |
252 |
< |
regions of the phase space. The condition of an ensemble at any time |
227 |
> |
|
228 |
> |
In statistical mechanics, the condition of an ensemble at any time |
229 |
|
can be regarded as appropriately specified by the density $\rho$ |
230 |
|
with which representative points are distributed over the phase |
231 |
|
space. The density distribution for an ensemble with $f$ degrees of |
280 |
|
statistical characteristics. As a function of macroscopic |
281 |
|
parameters, such as temperature \textit{etc}, the partition function |
282 |
|
can be used to describe the statistical properties of a system in |
283 |
< |
thermodynamic equilibrium. |
284 |
< |
|
285 |
< |
As an ensemble of systems, each of which is known to be thermally |
310 |
< |
isolated and conserve energy, the Microcanonical ensemble (NVE) has |
311 |
< |
a partition function like, |
283 |
> |
thermodynamic equilibrium. As an ensemble of systems, each of which |
284 |
> |
is known to be thermally isolated and conserve energy, the |
285 |
> |
Microcanonical ensemble (NVE) has a partition function like, |
286 |
|
\begin{equation} |
287 |
< |
\Omega (N,V,E) = e^{\beta TS} \label{introEquation:NVEPartition}. |
287 |
> |
\Omega (N,V,E) = e^{\beta TS}. \label{introEquation:NVEPartition}. |
288 |
|
\end{equation} |
289 |
|
A canonical ensemble (NVT)is an ensemble of systems, each of which |
290 |
|
can share its energy with a large heat reservoir. The distribution |
291 |
|
of the total energy amongst the possible dynamical states is given |
292 |
|
by the partition function, |
293 |
|
\begin{equation} |
294 |
< |
\Omega (N,V,T) = e^{ - \beta A} |
294 |
> |
\Omega (N,V,T) = e^{ - \beta A}. |
295 |
|
\label{introEquation:NVTPartition} |
296 |
|
\end{equation} |
297 |
|
Here, $A$ is the Helmholtz free energy which is defined as $ A = U - |
348 |
|
\frac{{\partial \rho }}{{\partial p_i }}\dot p_i } \right)} = 0 . |
349 |
|
\label{introEquation:liouvilleTheorem} |
350 |
|
\end{equation} |
377 |
– |
|
351 |
|
Liouville's theorem states that the distribution function is |
352 |
|
constant along any trajectory in phase space. In classical |
353 |
|
statistical mechanics, since the number of members in an ensemble is |
489 |
|
$\omega(\lambda_1x_1+\lambda_2x_2, y) = \lambda_1\omega(x_1, y)+ |
490 |
|
\lambda_2\omega(x_2, y)$, $\omega(x, y) = - \omega(y, x)$ and |
491 |
|
$\omega(x, x) = 0$. The cross product operation in vector field is |
492 |
< |
an example of symplectic form. |
493 |
< |
|
494 |
< |
One of the motivations to study \emph{symplectic manifolds} in |
495 |
< |
Hamiltonian Mechanics is that a symplectic manifold can represent |
496 |
< |
all possible configurations of the system and the phase space of the |
497 |
< |
system can be described by it's cotangent bundle. Every symplectic |
498 |
< |
manifold is even dimensional. For instance, in Hamilton equations, |
526 |
< |
coordinate and momentum always appear in pairs. |
492 |
> |
an example of symplectic form. One of the motivations to study |
493 |
> |
\emph{symplectic manifolds} in Hamiltonian Mechanics is that a |
494 |
> |
symplectic manifold can represent all possible configurations of the |
495 |
> |
system and the phase space of the system can be described by it's |
496 |
> |
cotangent bundle. Every symplectic manifold is even dimensional. For |
497 |
> |
instance, in Hamilton equations, coordinate and momentum always |
498 |
> |
appear in pairs. |
499 |
|
|
500 |
|
\subsection{\label{introSection:ODE}Ordinary Differential Equations} |
501 |
|
|
522 |
|
\frac{d}{{dt}}x = J\nabla _x H(x) |
523 |
|
\label{introEquation:compactHamiltonian} |
524 |
|
\end{equation}In this case, $f$ is |
525 |
< |
called a \emph{Hamiltonian vector field}. |
526 |
< |
|
555 |
< |
Another generalization of Hamiltonian dynamics is Poisson |
556 |
< |
Dynamics\cite{Olver1986}, |
525 |
> |
called a \emph{Hamiltonian vector field}. Another generalization of |
526 |
> |
Hamiltonian dynamics is Poisson Dynamics\cite{Olver1986}, |
527 |
|
\begin{equation} |
528 |
|
\dot x = J(x)\nabla _x H \label{introEquation:poissonHamiltonian} |
529 |
|
\end{equation} |
573 |
|
The hidden geometric properties\cite{Budd1999, Marsden1998} of an |
574 |
|
ODE and its flow play important roles in numerical studies. Many of |
575 |
|
them can be found in systems which occur naturally in applications. |
606 |
– |
|
576 |
|
Let $\varphi$ be the flow of Hamiltonian vector field, $\varphi$ is |
577 |
|
a \emph{symplectic} flow if it satisfies, |
578 |
|
\begin{equation} |
586 |
|
\begin{equation} |
587 |
|
{\varphi '}^T J \varphi ' = J \circ \varphi |
588 |
|
\end{equation} |
589 |
< |
is the property that must be preserved by the integrator. |
590 |
< |
|
591 |
< |
It is possible to construct a \emph{volume-preserving} flow for a |
592 |
< |
source free ODE ($ \nabla \cdot f = 0 $), if the flow satisfies $ |
593 |
< |
\det d\varphi = 1$. One can show easily that a symplectic flow will |
625 |
< |
be volume-preserving. |
626 |
< |
|
627 |
< |
Changing the variables $y = h(x)$ in an ODE |
589 |
> |
is the property that must be preserved by the integrator. It is |
590 |
> |
possible to construct a \emph{volume-preserving} flow for a source |
591 |
> |
free ODE ($ \nabla \cdot f = 0 $), if the flow satisfies $ \det |
592 |
> |
d\varphi = 1$. One can show easily that a symplectic flow will be |
593 |
> |
volume-preserving. Changing the variables $y = h(x)$ in an ODE |
594 |
|
(Eq.~\ref{introEquation:ODE}) will result in a new system, |
595 |
|
\[ |
596 |
|
\dot y = \tilde f(y) = ((dh \cdot f)h^{ - 1} )(y). |
597 |
|
\] |
598 |
|
The vector filed $f$ has reversing symmetry $h$ if $f = - \tilde f$. |
599 |
|
In other words, the flow of this vector field is reversible if and |
600 |
< |
only if $ h \circ \varphi ^{ - 1} = \varphi \circ h $. |
601 |
< |
|
636 |
< |
A \emph{first integral}, or conserved quantity of a general |
600 |
> |
only if $ h \circ \varphi ^{ - 1} = \varphi \circ h $. A |
601 |
> |
\emph{first integral}, or conserved quantity of a general |
602 |
|
differential function is a function $ G:R^{2d} \to R^d $ which is |
603 |
|
constant for all solutions of the ODE $\frac{{dx}}{{dt}} = f(x)$ , |
604 |
|
\[ |
611 |
|
which is the condition for conserving \emph{first integral}. For a |
612 |
|
canonical Hamiltonian system, the time evolution of an arbitrary |
613 |
|
smooth function $G$ is given by, |
649 |
– |
|
614 |
|
\begin{eqnarray} |
615 |
|
\frac{{dG(x(t))}}{{dt}} & = & [\nabla _x G(x(t))]^T \dot x(t) \\ |
616 |
|
& = & [\nabla _x G(x(t))]^T J\nabla _x H(x(t)). \\ |
617 |
|
\label{introEquation:firstIntegral1} |
618 |
|
\end{eqnarray} |
655 |
– |
|
656 |
– |
|
619 |
|
Using poisson bracket notion, Equation |
620 |
|
\ref{introEquation:firstIntegral1} can be rewritten as |
621 |
|
\[ |
628 |
|
\] |
629 |
|
As well known, the Hamiltonian (or energy) H of a Hamiltonian system |
630 |
|
is a \emph{first integral}, which is due to the fact $\{ H,H\} = |
631 |
< |
0$. |
670 |
< |
|
671 |
< |
When designing any numerical methods, one should always try to |
631 |
> |
0$. When designing any numerical methods, one should always try to |
632 |
|
preserve the structural properties of the original ODE and its flow. |
633 |
|
|
634 |
|
\subsection{\label{introSection:constructionSymplectic}Construction of Symplectic Methods} |
669 |
|
\label{introEquation:FlowDecomposition} |
670 |
|
\end{equation} |
671 |
|
where each of the sub-flow is chosen such that each represent a |
672 |
< |
simpler integration of the system. |
673 |
< |
|
714 |
< |
Suppose that a Hamiltonian system takes the form, |
672 |
> |
simpler integration of the system. Suppose that a Hamiltonian system |
673 |
> |
takes the form, |
674 |
|
\[ |
675 |
|
H = H_1 + H_2. |
676 |
|
\] |
711 |
|
\begin{equation} |
712 |
|
\varphi _h^{ - 1} = \varphi _{ - h}. |
713 |
|
\label{introEquation:timeReversible} |
714 |
< |
\end{equation},appendixFig:architecture |
714 |
> |
\end{equation} |
715 |
|
|
716 |
|
\subsubsection{\label{introSection:exampleSplittingMethod}\textbf{Examples of the Splitting Method}} |
717 |
|
The classical equation for a system consisting of interacting |
760 |
|
|
761 |
|
\item Repeat from step 1 with the new position, velocities, and forces assuming the roles of the initial values. |
762 |
|
\end{enumerate} |
804 |
– |
|
763 |
|
By simply switching the order of the propagators in the splitting |
764 |
|
and composing a new integrator, the \emph{position verlet} |
765 |
|
integrator, can be generated, |
1062 |
|
to justify the correctness of a liquid model. Moreover, various |
1063 |
|
equilibrium thermodynamic and structural properties can also be |
1064 |
|
expressed in terms of radial distribution function \cite{Allen1987}. |
1107 |
– |
|
1065 |
|
The pair distribution functions $g(r)$ gives the probability that a |
1066 |
|
particle $i$ will be located at a distance $r$ from a another |
1067 |
|
particle $j$ in the system |
1189 |
|
Q^T PJ^{ - 1} + J^{ - 1} P^T Q = 0 . \\ |
1190 |
|
\label{introEquation:RBFirstOrderConstraint} |
1191 |
|
\end{equation} |
1235 |
– |
|
1192 |
|
Using Equation (\ref{introEquation:motionHamiltonianCoordinate}, |
1193 |
|
\ref{introEquation:motionHamiltonianMomentum}), one can write down |
1194 |
|
the equations of motion, |
1239 |
– |
|
1195 |
|
\begin{eqnarray} |
1196 |
|
\frac{{dq}}{{dt}} & = & \frac{p}{m} \label{introEquation:RBMotionPosition}\\ |
1197 |
|
\frac{{dp}}{{dt}} & = & - \nabla _q V(q,Q) \label{introEquation:RBMotionMomentum}\\ |
1198 |
|
\frac{{dQ}}{{dt}} & = & PJ^{ - 1} \label{introEquation:RBMotionRotation}\\ |
1199 |
|
\frac{{dP}}{{dt}} & = & - \nabla _Q V(q,Q) - 2Q\Lambda . \label{introEquation:RBMotionP} |
1200 |
|
\end{eqnarray} |
1246 |
– |
|
1201 |
|
In general, there are two ways to satisfy the holonomic constraints. |
1202 |
|
We can use a constraint force provided by a Lagrange multiplier on |
1203 |
|
the normal manifold to keep the motion on constraint space. Or we |
1209 |
|
M = \left\{ {(Q,P):Q^T Q = 1,Q^T PJ^{ - 1} + J^{ - 1} P^T Q = 0} |
1210 |
|
\right\}. |
1211 |
|
\] |
1258 |
– |
|
1212 |
|
Unfortunately, this constraint manifold is not the cotangent bundle |
1213 |
< |
$T_{\star}SO(3)$. However, it turns out that under symplectic |
1213 |
> |
$T^* SO(3)$ which can be consider as a symplectic manifold on Lie |
1214 |
> |
rotation group $SO(3)$. However, it turns out that under symplectic |
1215 |
|
transformation, the cotangent space and the phase space are |
1216 |
|
diffeomorphic. By introducing |
1217 |
|
\[ |
1223 |
|
T^* SO(3) = \left\{ {(\tilde Q,\tilde P):\tilde Q^T \tilde Q = |
1224 |
|
1,\tilde Q^T \tilde PJ^{ - 1} + J^{ - 1} P^T \tilde Q = 0} \right\} |
1225 |
|
\] |
1272 |
– |
|
1226 |
|
For a body fixed vector $X_i$ with respect to the center of mass of |
1227 |
|
the rigid body, its corresponding lab fixed vector $X_0^{lab}$ is |
1228 |
|
given as |
1241 |
|
\[ |
1242 |
|
\nabla _Q V(q,Q) = F(q,Q)X_i^t |
1243 |
|
\] |
1244 |
< |
respectively. |
1245 |
< |
|
1246 |
< |
As a common choice to describe the rotation dynamics of the rigid |
1294 |
< |
body, the angular momentum on the body fixed frame $\Pi = Q^t P$ is |
1295 |
< |
introduced to rewrite the equations of motion, |
1244 |
> |
respectively. As a common choice to describe the rotation dynamics |
1245 |
> |
of the rigid body, the angular momentum on the body fixed frame $\Pi |
1246 |
> |
= Q^t P$ is introduced to rewrite the equations of motion, |
1247 |
|
\begin{equation} |
1248 |
|
\begin{array}{l} |
1249 |
< |
\mathop \Pi \limits^ \bullet = J^{ - 1} \Pi ^T \Pi + Q^T \sum\limits_i {F_i (q,Q)X_i^T } - \Lambda \\ |
1250 |
< |
\mathop Q\limits^{{\rm{ }} \bullet } = Q\Pi {\rm{ }}J^{ - 1} \\ |
1249 |
> |
\dot \Pi = J^{ - 1} \Pi ^T \Pi + Q^T \sum\limits_i {F_i (q,Q)X_i^T } - \Lambda, \\ |
1250 |
> |
\dot Q = Q\Pi {\rm{ }}J^{ - 1}, \\ |
1251 |
|
\end{array} |
1252 |
|
\label{introEqaution:RBMotionPI} |
1253 |
|
\end{equation} |
1254 |
< |
, as well as holonomic constraints, |
1254 |
> |
as well as holonomic constraints, |
1255 |
|
\[ |
1256 |
|
\begin{array}{l} |
1257 |
< |
\Pi J^{ - 1} + J^{ - 1} \Pi ^t = 0 \\ |
1258 |
< |
Q^T Q = 1 \\ |
1257 |
> |
\Pi J^{ - 1} + J^{ - 1} \Pi ^t = 0, \\ |
1258 |
> |
Q^T Q = 1 .\\ |
1259 |
|
\end{array} |
1260 |
|
\] |
1310 |
– |
|
1261 |
|
For a vector $v(v_1 ,v_2 ,v_3 ) \in R^3$ and a matrix $\hat v \in |
1262 |
|
so(3)^ \star$, the hat-map isomorphism, |
1263 |
|
\begin{equation} |
1272 |
|
will let us associate the matrix products with traditional vector |
1273 |
|
operations |
1274 |
|
\[ |
1275 |
< |
\hat vu = v \times u |
1275 |
> |
\hat vu = v \times u. |
1276 |
|
\] |
1277 |
< |
Using \ref{introEqaution:RBMotionPI}, one can construct a skew |
1277 |
> |
Using Eq.~\ref{introEqaution:RBMotionPI}, one can construct a skew |
1278 |
|
matrix, |
1279 |
+ |
\begin{eqnarray} |
1280 |
+ |
(\dot \Pi - \dot \Pi ^T ){\rm{ }} &= &{\rm{ }}(\Pi - \Pi ^T ){\rm{ |
1281 |
+ |
}}(J^{ - 1} \Pi + \Pi J^{ - 1} ) \notag \\ |
1282 |
+ |
+ \sum\limits_i {[Q^T F_i |
1283 |
+ |
(r,Q)X_i^T - X_i F_i (r,Q)^T Q]} - (\Lambda - \Lambda ^T ). |
1284 |
+ |
\label{introEquation:skewMatrixPI} |
1285 |
+ |
\end{eqnarray} |
1286 |
+ |
Since $\Lambda$ is symmetric, the last term of |
1287 |
+ |
Eq.~\ref{introEquation:skewMatrixPI} is zero, which implies the |
1288 |
+ |
Lagrange multiplier $\Lambda$ is absent from the equations of |
1289 |
+ |
motion. This unique property eliminates the requirement of |
1290 |
+ |
iterations which can not be avoided in other methods\cite{Kol1997, |
1291 |
+ |
Omelyan1998}. Applying the hat-map isomorphism, we obtain the |
1292 |
+ |
equation of motion for angular momentum on body frame |
1293 |
|
\begin{equation} |
1330 |
– |
(\mathop \Pi \limits^ \bullet - \mathop \Pi \limits^ {\bullet ^T} |
1331 |
– |
){\rm{ }} = {\rm{ }}(\Pi - \Pi ^T ){\rm{ }}(J^{ - 1} \Pi + \Pi J^{ |
1332 |
– |
- 1} ) + \sum\limits_i {[Q^T F_i (r,Q)X_i^T - X_i F_i (r,Q)^T Q]} - |
1333 |
– |
(\Lambda - \Lambda ^T ) . \label{introEquation:skewMatrixPI} |
1334 |
– |
\end{equation} |
1335 |
– |
Since $\Lambda$ is symmetric, the last term of Equation |
1336 |
– |
\ref{introEquation:skewMatrixPI} is zero, which implies the Lagrange |
1337 |
– |
multiplier $\Lambda$ is absent from the equations of motion. This |
1338 |
– |
unique property eliminates the requirement of iterations which can |
1339 |
– |
not be avoided in other methods\cite{Kol1997, Omelyan1998}. |
1340 |
– |
|
1341 |
– |
Applying the hat-map isomorphism, we obtain the equation of motion |
1342 |
– |
for angular momentum on body frame |
1343 |
– |
\begin{equation} |
1294 |
|
\dot \pi = \pi \times I^{ - 1} \pi + \sum\limits_i {\left( {Q^T |
1295 |
|
F_i (r,Q)} \right) \times X_i }. |
1296 |
|
\label{introEquation:bodyAngularMotion} |
1298 |
|
In the same manner, the equation of motion for rotation matrix is |
1299 |
|
given by |
1300 |
|
\[ |
1301 |
< |
\dot Q = Qskew(I^{ - 1} \pi ) |
1301 |
> |
\dot Q = Qskew(I^{ - 1} \pi ). |
1302 |
|
\] |
1303 |
|
|
1304 |
|
\subsection{\label{introSection:SymplecticFreeRB}Symplectic |
1320 |
|
0 & {\pi _3 } & { - \pi _2 } \\ |
1321 |
|
{ - \pi _3 } & 0 & {\pi _1 } \\ |
1322 |
|
{\pi _2 } & { - \pi _1 } & 0 \\ |
1323 |
< |
\end{array}} \right) |
1323 |
> |
\end{array}} \right). |
1324 |
|
\end{equation} |
1325 |
|
Thus, the dynamics of free rigid body is governed by |
1326 |
|
\begin{equation} |
1327 |
< |
\frac{d}{{dt}}\pi = J(\pi )\nabla _\pi T^r (\pi ) |
1327 |
> |
\frac{d}{{dt}}\pi = J(\pi )\nabla _\pi T^r (\pi ). |
1328 |
|
\end{equation} |
1379 |
– |
|
1329 |
|
One may notice that each $T_i^r$ in Equation |
1330 |
|
\ref{introEquation:rotationalKineticRB} can be solved exactly. For |
1331 |
|
instance, the equations of motion due to $T_1^r$ are given by |
1358 |
|
propagator, |
1359 |
|
\[ |
1360 |
|
e^{\Delta tR_1 } \approx (1 - \Delta tR_1 )^{ - 1} (1 + \Delta tR_1 |
1361 |
< |
) |
1361 |
> |
). |
1362 |
|
\] |
1363 |
|
The flow maps for $T_2^r$ and $T_3^r$ can be found in the same |
1364 |
|
manner. In order to construct a second-order symplectic method, we |
1376 |
|
\circ \varphi _{\Delta t/2,\pi _2 } \circ \varphi _{\Delta t/2,\pi |
1377 |
|
_1 }. |
1378 |
|
\] |
1430 |
– |
|
1379 |
|
The non-canonical Lie-Poisson bracket ${F, G}$ of two function |
1380 |
|
$F(\pi )$ and $G(\pi )$ is defined by |
1381 |
|
\[ |
1382 |
|
\{ F,G\} (\pi ) = [\nabla _\pi F(\pi )]^T J(\pi )\nabla _\pi G(\pi |
1383 |
< |
) |
1383 |
> |
). |
1384 |
|
\] |
1385 |
|
If the Poisson bracket of a function $F$ with an arbitrary smooth |
1386 |
|
function $G$ is zero, $F$ is a \emph{Casimir}, which is the |
1391 |
|
then by the chain rule |
1392 |
|
\[ |
1393 |
|
\nabla _\pi F(\pi ) = S'(\frac{{\parallel \pi \parallel ^2 |
1394 |
< |
}}{2})\pi |
1394 |
> |
}}{2})\pi. |
1395 |
|
\] |
1396 |
< |
Thus $ [\nabla _\pi F(\pi )]^T J(\pi ) = - S'(\frac{{\parallel \pi |
1396 |
> |
Thus, $ [\nabla _\pi F(\pi )]^T J(\pi ) = - S'(\frac{{\parallel |
1397 |
> |
\pi |
1398 |
|
\parallel ^2 }}{2})\pi \times \pi = 0 $. This explicit |
1399 |
|
Lie-Poisson integrator is found to be both extremely efficient and |
1400 |
|
stable. These properties can be explained by the fact the small |
1407 |
|
The Hamiltonian of rigid body can be separated in terms of kinetic |
1408 |
|
energy and potential energy, |
1409 |
|
\[ |
1410 |
< |
H = T(p,\pi ) + V(q,Q) |
1410 |
> |
H = T(p,\pi ) + V(q,Q). |
1411 |
|
\] |
1412 |
|
The equations of motion corresponding to potential energy and |
1413 |
|
kinetic energy are listed in the below table, |
1414 |
|
\begin{table} |
1415 |
< |
\caption{Equations of motion due to Potential and Kinetic Energies} |
1415 |
> |
\caption{EQUATIONS OF MOTION DUE TO POTENTIAL AND KINETIC ENERGIES} |
1416 |
|
\begin{center} |
1417 |
|
\begin{tabular}{|l|l|} |
1418 |
|
\hline |
1456 |
|
\] |
1457 |
|
Finally, we obtain the overall symplectic propagators for freely |
1458 |
|
moving rigid bodies |
1459 |
< |
\begin{equation} |
1460 |
< |
\begin{array}{c} |
1461 |
< |
\varphi _{\Delta t} = \varphi _{\Delta t/2,F} \circ \varphi _{\Delta t/2,\tau } \\ |
1462 |
< |
\circ \varphi _{\Delta t,T^t } \circ \varphi _{\Delta t/2,\pi _1 } \circ \varphi _{\Delta t/2,\pi _2 } \circ \varphi _{\Delta t,\pi _3 } \circ \varphi _{\Delta t/2,\pi _2 } \circ \varphi _{\Delta t/2,\pi _1 } \\ |
1514 |
< |
\circ \varphi _{\Delta t/2,\tau } \circ \varphi _{\Delta t/2,F} .\\ |
1515 |
< |
\end{array} |
1459 |
> |
\begin{eqnarray*} |
1460 |
> |
\varphi _{\Delta t} &=& \varphi _{\Delta t/2,F} \circ \varphi _{\Delta t/2,\tau } \\ |
1461 |
> |
& & \circ \varphi _{\Delta t,T^t } \circ \varphi _{\Delta t/2,\pi _1 } \circ \varphi _{\Delta t/2,\pi _2 } \circ \varphi _{\Delta t,\pi _3 } \circ \varphi _{\Delta t/2,\pi _2 } \circ \varphi _{\Delta t/2,\pi _1 } \\ |
1462 |
> |
& & \circ \varphi _{\Delta t/2,\tau } \circ \varphi _{\Delta t/2,F} .\\ |
1463 |
|
\label{introEquation:overallRBFlowMaps} |
1464 |
< |
\end{equation} |
1464 |
> |
\end{eqnarray*} |
1465 |
|
|
1466 |
|
\section{\label{introSection:langevinDynamics}Langevin Dynamics} |
1467 |
|
As an alternative to newtonian dynamics, Langevin dynamics, which |
1507 |
|
\[ |
1508 |
|
W(x) = U(x) - \sum\limits_{\alpha = 1}^N {\frac{{g_\alpha ^2 |
1509 |
|
}}{{2m_\alpha w_\alpha ^2 }}} x^2 |
1510 |
< |
\] and combining the last two terms in Equation |
1511 |
< |
\ref{introEquation:bathGLE}, we may rewrite the Harmonic bath |
1565 |
< |
Hamiltonian as |
1510 |
> |
\] |
1511 |
> |
and combining the last two terms in Eq.~\ref{introEquation:bathGLE}, we may rewrite the Harmonic bath Hamiltonian as |
1512 |
|
\[ |
1513 |
|
H = \frac{{p^2 }}{{2m}} + W(x) + \sum\limits_{\alpha = 1}^N |
1514 |
|
{\left\{ {\frac{{p_\alpha ^2 }}{{2m_\alpha }} + \frac{1}{2}m_\alpha |
1515 |
|
w_\alpha ^2 \left( {x_\alpha - \frac{{g_\alpha }}{{m_\alpha |
1516 |
< |
w_\alpha ^2 }}x} \right)^2 } \right\}} |
1516 |
> |
w_\alpha ^2 }}x} \right)^2 } \right\}}. |
1517 |
|
\] |
1518 |
|
Since the first two terms of the new Hamiltonian depend only on the |
1519 |
|
system coordinates, we can get the equations of motion for |
1530 |
|
\frac{{g_\alpha }}{{m_\alpha w_\alpha ^2 }}x} \right). |
1531 |
|
\label{introEquation:bathMotionGLE} |
1532 |
|
\end{equation} |
1587 |
– |
|
1533 |
|
In order to derive an equation for $x$, the dynamics of the bath |
1534 |
|
variables $x_\alpha$ must be solved exactly first. As an integral |
1535 |
|
transform which is particularly useful in solving linear ordinary |
1538 |
|
differential equations into simple algebra problems which can be |
1539 |
|
solved easily. Then, by applying the inverse Laplace transform, also |
1540 |
|
known as the Bromwich integral, we can retrieve the solutions of the |
1541 |
< |
original problems. |
1542 |
< |
|
1598 |
< |
Let $f(t)$ be a function defined on $ [0,\infty ) $. The Laplace |
1599 |
< |
transform of f(t) is a new function defined as |
1541 |
> |
original problems. Let $f(t)$ be a function defined on $ [0,\infty ) |
1542 |
> |
$. The Laplace transform of f(t) is a new function defined as |
1543 |
|
\[ |
1544 |
|
L(f(t)) \equiv F(p) = \int_0^\infty {f(t)e^{ - pt} dt} |
1545 |
|
\] |
1546 |
|
where $p$ is real and $L$ is called the Laplace Transform |
1547 |
|
Operator. Below are some important properties of Laplace transform |
1605 |
– |
|
1548 |
|
\begin{eqnarray*} |
1549 |
|
L(x + y) & = & L(x) + L(y) \\ |
1550 |
|
L(ax) & = & aL(x) \\ |
1552 |
|
L(\ddot x)& = & p^2 L(x) - px(0) - \dot x(0) \\ |
1553 |
|
L\left( {\int_0^t {g(t - \tau )h(\tau )d\tau } } \right)& = & G(p)H(p) \\ |
1554 |
|
\end{eqnarray*} |
1613 |
– |
|
1614 |
– |
|
1555 |
|
Applying the Laplace transform to the bath coordinates, we obtain |
1556 |
|
\begin{eqnarray*} |
1557 |
|
p^2 L(x_\alpha ) - px_\alpha (0) - \dot x_\alpha (0) & = & - \omega _\alpha ^2 L(x_\alpha ) + \frac{{g_\alpha }}{{\omega _\alpha }}L(x) \\ |
1558 |
|
L(x_\alpha ) & = & \frac{{\frac{{g_\alpha }}{{\omega _\alpha }}L(x) + px_\alpha (0) + \dot x_\alpha (0)}}{{p^2 + \omega _\alpha ^2 }} \\ |
1559 |
|
\end{eqnarray*} |
1620 |
– |
|
1560 |
|
By the same way, the system coordinates become |
1561 |
|
\begin{eqnarray*} |
1562 |
< |
mL(\ddot x) & = & - \frac{1}{p}\frac{{\partial W(x)}}{{\partial x}} \\ |
1563 |
< |
& & \mbox{} - \sum\limits_{\alpha = 1}^N {\left\{ { - \frac{{g_\alpha ^2 }}{{m_\alpha \omega _\alpha ^2 }}\frac{p}{{p^2 + \omega _\alpha ^2 }}pL(x) - \frac{p}{{p^2 + \omega _\alpha ^2 }}g_\alpha x_\alpha (0) - \frac{1}{{p^2 + \omega _\alpha ^2 }}g_\alpha \dot x_\alpha (0)} \right\}} \\ |
1562 |
> |
mL(\ddot x) & = & |
1563 |
> |
- \sum\limits_{\alpha = 1}^N {\left\{ { - \frac{{g_\alpha ^2 }}{{m_\alpha \omega _\alpha ^2 }}\frac{p}{{p^2 + \omega _\alpha ^2 }}pL(x) - \frac{p}{{p^2 + \omega _\alpha ^2 }}g_\alpha x_\alpha (0) - \frac{1}{{p^2 + \omega _\alpha ^2 }}g_\alpha \dot x_\alpha (0)} \right\}} \\ |
1564 |
> |
& & - \frac{1}{p}\frac{{\partial W(x)}}{{\partial x}} |
1565 |
|
\end{eqnarray*} |
1626 |
– |
|
1566 |
|
With the help of some relatively important inverse Laplace |
1567 |
|
transformations: |
1568 |
|
\[ |
1572 |
|
L(1) = \frac{1}{p} \\ |
1573 |
|
\end{array} |
1574 |
|
\] |
1575 |
< |
, we obtain |
1575 |
> |
we obtain |
1576 |
|
\begin{eqnarray*} |
1577 |
|
m\ddot x & = & - \frac{{\partial W(x)}}{{\partial x}} - |
1578 |
|
\sum\limits_{\alpha = 1}^N {\left\{ {\left( { - \frac{{g_\alpha ^2 |
1645 |
|
\[ |
1646 |
|
\int_0^t {\xi (t)\dot x(t - \tau )d\tau } = \xi _0 (x(t) - x(0)) |
1647 |
|
\] |
1648 |
< |
and Equation \ref{introEuqation:GeneralizedLangevinDynamics} becomes |
1648 |
> |
and Eq.~\ref{introEuqation:GeneralizedLangevinDynamics} becomes |
1649 |
|
\[ |
1650 |
|
m\ddot x = - \frac{\partial }{{\partial x}}\left( {W(x) + |
1651 |
|
\frac{1}{2}\xi _0 (x - x_0 )^2 } \right) + R(t), |
1662 |
|
\int_0^t {\xi (t)\dot x(t - \tau )d\tau } = 2\xi _0 \int_0^t |
1663 |
|
{\delta (t)\dot x(t - \tau )d\tau } = \xi _0 \dot x(t), |
1664 |
|
\] |
1665 |
< |
and Equation \ref{introEuqation:GeneralizedLangevinDynamics} becomes |
1665 |
> |
and Eq.~\ref{introEuqation:GeneralizedLangevinDynamics} becomes |
1666 |
|
\begin{equation} |
1667 |
|
m\ddot x = - \frac{{\partial W(x)}}{{\partial x}} - \xi _0 \dot |
1668 |
|
x(t) + R(t) \label{introEquation:LangevinEquation} |
1685 |
|
R(t) = \sum\limits_{\alpha = 1}^N {g_\alpha q_\alpha (t)}. |
1686 |
|
\] |
1687 |
|
And since the $q$ coordinates are harmonic oscillators, |
1749 |
– |
|
1688 |
|
\begin{eqnarray*} |
1689 |
|
\left\langle {q_\alpha ^2 } \right\rangle & = & \frac{{kT}}{{m_\alpha \omega _\alpha ^2 }} \\ |
1690 |
|
\left\langle {q_\alpha (t)q_\alpha (0)} \right\rangle & = & \left\langle {q_\alpha ^2 (0)} \right\rangle \cos (\omega _\alpha t) \\ |
1693 |
|
& = &\sum\limits_\alpha {g_\alpha ^2 \left\langle {q_\alpha ^2 (0)} \right\rangle \cos (\omega _\alpha t)} \\ |
1694 |
|
& = &kT\xi (t) \\ |
1695 |
|
\end{eqnarray*} |
1758 |
– |
|
1696 |
|
Thus, we recover the \emph{second fluctuation dissipation theorem} |
1697 |
|
\begin{equation} |
1698 |
|
\xi (t) = \left\langle {R(t)R(0)} \right\rangle |