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\begin{document} |
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|
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\title{Langevin Dynamics for Rigid Body of Arbitrary Shape } |
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\title{An algorithm for performing Langevin dynamics on rigid bodies of arbitrary shape } |
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|
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\author{Teng Lin, Xiuquan Sun and J. Daniel Gezelter\footnote{Corresponding author. \ Electronic mail: |
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\author{Xiuquan Sun, Teng Lin and J. Daniel Gezelter\footnote{Corresponding author. \ Electronic mail: |
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gezelter@nd.edu} \\ |
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Department of Chemistry and Biochemistry\\ |
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University of Notre Dame\\ |
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algorithm for arbitrary-shaped rigid particles by integrating the |
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accurate estimation of friction tensor from hydrodynamics theory |
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into the sophisticated rigid body dynamics algorithms. |
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|
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\section{Computational Methods{\label{methodSec}}} |
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|
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\subsection{\label{introSection:frictionTensor}Friction Tensor} |
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Theoretically, the friction kernel can be determined using the |
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\begin{eqnarray} |
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\Xi _{}^{tt} & = & \sum\limits_i {\sum\limits_j {C_{ij} } } \notag , \\ |
| 367 |
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\Xi _{}^{tr} & = & \Xi _{}^{rt} = \sum\limits_i {\sum\limits_j {U_i C_{ij} } } , \\ |
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\Xi _{}^{rr} & = & - \sum\limits_i {\sum\limits_j {U_i C_{ij} } } U_j. \notag \\ |
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\Xi _{}^{rr} & = & - \sum\limits_i {\sum\limits_j {U_i C_{ij} } } |
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U_j + 6 \eta V {\bf I}. \notag |
| 370 |
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\label{introEquation:ResistanceTensorArbitraryOrigin} |
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\end{eqnarray} |
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The final term in the expression for $\Xi^{rr}$ is correction that |
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+ |
accounts for errors in the rotational motion of certain kinds of bead |
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+ |
models. The additive correction uses the solvent viscosity ($\eta$) |
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+ |
as well as the total volume of the beads that contribute to the |
| 376 |
+ |
hydrodynamic model, |
| 377 |
+ |
\begin{equation} |
| 378 |
+ |
V = \frac{4 \pi}{3} \sum_{i=1}^{N} \sigma_i^3, |
| 379 |
+ |
\end{equation} |
| 380 |
+ |
where $\sigma_i$ is the radius of bead $i$. This correction term was |
| 381 |
+ |
rigorously tested and compared with the analytical results for |
| 382 |
+ |
two-sphere and ellipsoidal systems by Garcia de la Torre and |
| 383 |
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Rodes.\cite{Torre:1983lr} |
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+ |
|
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+ |
|
| 386 |
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The resistance tensor depends on the origin to which they refer. The |
| 387 |
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proper location for applying the friction force is the center of |
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resistance (or center of reaction), at which the trace of rotational |
| 439 |
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where $x_OR$, $y_OR$, $z_OR$ are the components of the vector |
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joining center of resistance $R$ and origin $O$. |
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|
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\subsection{Langevin Dynamics for Rigid Particles of Arbitrary Shape\label{LDRB}} |
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|
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\section{Langevin Dynamics for Rigid Particles of Arbitrary Shape\label{LDRB}} |
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Consider the Langevin equations of motion in generalized coordinates |
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|
\begin{equation} |
| 446 |
|
M_i \dot V_i (t) = F_{s,i} (t) + F_{f,i(t)} + F_{r,i} (t) |
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+ \frac{h}{2} {\bf \tau}^b(t + h) . |
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\end{align*} |
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|
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\section{Results and Discussion} |
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|
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The Langevin algorithm described in previous section has been |
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implemented in {\sc oopse}\cite{Meineke2005} and applied to studies |
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of the static and dynamic properties in several systems. |
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|
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\subsection{Temperature Control} |
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\section{Validating the Method\label{sec:validating}} |
| 591 |
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In order to validate our Langevin integrator for arbitrarily-shaped |
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rigid bodies, we implemented the algorithm in {\sc |
| 593 |
> |
oopse}\cite{Meineke2005} and compared the results of this algorithm |
| 594 |
> |
with the known |
| 595 |
> |
hydrodynamic limiting behavior for a few model systems, and to |
| 596 |
> |
microcanonical molecular dynamics simulations for some more |
| 597 |
> |
complicated bodies. The model systems and their analytical behavior |
| 598 |
> |
(if known) are summarized below. Parameters for the primary particles |
| 599 |
> |
comprising our model systems are given in table \ref{tab:parameters}, |
| 600 |
> |
and a sketch of the arrangement of these primary particles into the |
| 601 |
> |
model rigid bodies is shown in figure \ref{fig:models}. In table |
| 602 |
> |
\ref{tab:parameters}, $d$ and $l$ are the physical dimensions of |
| 603 |
> |
ellipsoidal (Gay-Berne) particles. For spherical particles, the value |
| 604 |
> |
of the Lennard-Jones $\sigma$ parameter is the particle diameter |
| 605 |
> |
($d$). Gay-Berne ellipsoids have an energy scaling parameter, |
| 606 |
> |
$\epsilon^s$, which describes the well depth for two identical |
| 607 |
> |
ellipsoids in a {\it side-by-side} configuration. Additionally, a |
| 608 |
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well depth aspect ratio, $\epsilon^r = \epsilon^e / \epsilon^s$, |
| 609 |
> |
describes the ratio between the well depths in the {\it end-to-end} |
| 610 |
> |
and side-by-side configurations. For spheres, $\epsilon^r \equiv 1$. |
| 611 |
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Moments of inertia are also required to describe the motion of primary |
| 612 |
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particles with orientational degrees of freedom. |
| 613 |
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|
| 614 |
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As shown in Eq.~\ref{randomForce}, random collisions associated with |
| 615 |
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the solvent's thermal motions is controlled by the external |
| 588 |
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temperature. The capability to maintain the temperature of the whole |
| 589 |
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system was usually used to measure the stability and efficiency of |
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the algorithm. In order to verify the stability of this new |
| 591 |
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algorithm, a series of simulations are performed on system |
| 592 |
< |
consisiting of 256 SSD water molecules with different viscosities. |
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The initial configuration for the simulations is taken from a 1ns |
| 594 |
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NVT simulation with a cubic box of 19.7166~\AA. All simulation are |
| 595 |
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carried out with cutoff radius of 9~\AA and 2 fs time step for 1 ns |
| 596 |
< |
with reference temperature at 300~K. The average temperature as a |
| 597 |
< |
function of $\eta$ is shown in Table \ref{langevin:viscosity} where |
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the temperatures range from 303.04~K to 300.47~K for $\eta = 0.01 - |
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< |
1$ poise. The better temperature control at higher viscosity can be |
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< |
explained by the finite size effect and relative slow relaxation |
| 601 |
< |
rate at lower viscosity regime. |
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< |
\begin{table} |
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\caption{AVERAGE TEMPERATURES FROM LANGEVIN DYNAMICS SIMULATIONS OF |
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SSD WATER MOLECULES WITH REFERENCE TEMPERATURE AT 300~K.} |
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\label{langevin:viscosity} |
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\begin{table*} |
| 615 |
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\begin{minipage}{\linewidth} |
| 616 |
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\begin{center} |
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< |
\begin{tabular}{lll} |
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\hline |
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$\eta$ & $\text{T}_{\text{avg}}$ & $\text{T}_{\text{rms}}$ \\ |
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\hline |
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1 & 300.47 & 10.99 \\ |
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0.1 & 301.19 & 11.136 \\ |
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0.01 & 303.04 & 11.796 \\ |
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\hline |
| 617 |
> |
\caption{Parameters for the primary particles in use by the rigid body |
| 618 |
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models in figure \ref{fig:models}.} |
| 619 |
> |
\begin{tabular}{lrcccccccc} |
| 620 |
> |
\hline |
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& & & & & & & \multicolumn{3}c{$\overleftrightarrow{\mathsf I}$ (amu \AA$^2$)} \\ |
| 622 |
> |
& & $d$ (\AA) & $l$ (\AA) & $\epsilon^s$ (kcal/mol) & $\epsilon^r$ & |
| 623 |
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$m$ (amu) & $I_{xx}$ & $I_{yy}$ & $I_{zz}$ \\ \hline |
| 624 |
> |
Sphere & & 6.5 & $= d$ & 0.8 & 1 & 190 & 802.75 & 802.75 & 802.75 \\ |
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Ellipsoid & & 4.6 & 13.8 & 0.8 & 0.2 & 200 & 2105 & 2105 & 421 \\ |
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> |
Dumbbell &(2 identical spheres) & 6.5 & $= d$ & 0.8 & 1 & 190 & 802.75 & 802.75 & 802.75 \\ |
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> |
Banana &(3 identical ellipsoids)& 4.2 & 11.2 & 0.8 & 0.2 & 240 & 10000 & 10000 & 0 \\ |
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> |
Lipid: & Spherical Head & 6.5 & $= d$ & 0.185 & 1 & 196 & & & \\ |
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> |
& Ellipsoidal Tail & 4.6 & 13.8 & 0.8 & 0.2 & 760 & 45000 & 45000 & 9000 \\ |
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> |
Solvent & & 4.7 & $= d$ & 0.8 & 1 & 72.06 & & & \\ |
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> |
\hline |
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|
\end{tabular} |
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\label{tab:parameters} |
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|
\end{center} |
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< |
\end{table} |
| 635 |
> |
\end{minipage} |
| 636 |
> |
\end{table*} |
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|
|
| 619 |
– |
Another set of calculations were performed to study the efficiency of |
| 620 |
– |
temperature control using different temperature coupling schemes. |
| 621 |
– |
The starting configuration is cooled to 173~K and evolved using NVE, |
| 622 |
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NVT, and Langevin dynamic with time step of 2 fs. |
| 623 |
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Fig.~\ref{langevin:temperature} shows the heating curve obtained as |
| 624 |
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the systems reach equilibrium. The orange curve in |
| 625 |
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Fig.~\ref{langevin:temperature} represents the simulation using |
| 626 |
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Nos\'e-Hoover temperature scaling scheme with thermostat of 5 ps |
| 627 |
– |
which gives reasonable tight coupling, while the blue one from |
| 628 |
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Langevin dynamics with viscosity of 0.1 poise demonstrates a faster |
| 629 |
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scaling to the desire temperature. When $ \eta = 0$, Langevin dynamics becomes normal |
| 630 |
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NVE (see orange curve in Fig.~\ref{langevin:temperature}) which |
| 631 |
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loses the temperature control ability. |
| 632 |
– |
|
| 638 |
|
\begin{figure} |
| 639 |
|
\centering |
| 640 |
< |
\includegraphics[width=\linewidth]{temperature.eps} |
| 641 |
< |
\caption[Plot of Temperature Fluctuation Versus Time]{Plot of |
| 642 |
< |
temperature fluctuation versus time.} \label{langevin:temperature} |
| 640 |
> |
\includegraphics[width=3in]{sketch} |
| 641 |
> |
\caption[Sketch of the model systems]{A sketch of the model systems |
| 642 |
> |
used in evaluating the behavior of the rigid body Langevin |
| 643 |
> |
integrator.} \label{fig:models} |
| 644 |
|
\end{figure} |
| 645 |
|
|
| 646 |
< |
\subsection{Langevin Dynamics of Banana Shaped Molecules} |
| 646 |
> |
\subsection{Simulation Methodology} |
| 647 |
> |
We performed reference microcanonical simulations with explicit |
| 648 |
> |
solvents for each of the different model system. In each case there |
| 649 |
> |
was one solute model and 1929 solvent molecules present in the |
| 650 |
> |
simulation box. All simulations were equilibrated using a |
| 651 |
> |
constant-pressure and temperature integrator with target values of 300 |
| 652 |
> |
K for the temperature and 1 atm for pressure. Following this stage, |
| 653 |
> |
further equilibration and sampling was done in a microcanonical |
| 654 |
> |
ensemble. Since the model bodies are typically quite massive, we were |
| 655 |
> |
able to use a time step of 25 fs. |
| 656 |
|
|
| 657 |
< |
In order to verify that Langevin dynamics can mimic the dynamics of |
| 658 |
< |
the systems absent of explicit solvents, we carried out two sets of |
| 659 |
< |
simulations and compare their dynamic properties. |
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< |
Fig.~\ref{langevin:twoBanana} shows a snapshot of the simulation |
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< |
made of 256 pentane molecules and two banana shaped molecules at |
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< |
273~K. It has an equivalent implicit solvent system containing only |
| 663 |
< |
two banana shaped molecules with viscosity of 0.289 center poise. To |
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< |
calculate the hydrodynamic properties of the banana shaped molecule, |
| 665 |
< |
we created a rough shell model (see Fig.~\ref{langevin:roughShell}), |
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< |
in which the banana shaped molecule is represented as a ``shell'' |
| 667 |
< |
made of 2266 small identical beads with size of 0.3 \AA on the |
| 668 |
< |
surface. Applying the procedure described in |
| 669 |
< |
Sec.~\ref{introEquation:ResistanceTensorArbitraryOrigin}, we |
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< |
identified the center of resistance at (0 $\rm{\AA}$, 0.7482 $\rm{\AA}$, |
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< |
-0.1988 $\rm{\AA}$), as well as the resistance tensor, |
| 672 |
< |
\[ |
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< |
\left( {\begin{array}{*{20}c} |
| 674 |
< |
0.9261 & 0 & 0&0&0.08585&0.2057\\ |
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< |
0& 0.9270&-0.007063& 0.08585&0&0\\ |
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< |
0&-0.007063&0.7494&0.2057&0&0\\ |
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< |
0&0.0858&0.2057& 58.64& 0&0\\ |
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< |
0.08585&0&0&0&48.30&3.219&\\ |
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< |
0.2057&0&0&0&3.219&10.7373\\ |
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< |
\end{array}} \right). |
| 666 |
< |
\] |
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< |
where the units for translational, translation-rotation coupling and rotational tensors are $\frac{kcal \cdot fs}{mol \cdot \rm{\AA}^2}$, $\frac{kcal \cdot fs}{mol \cdot \rm{\AA} \cdot rad}$ and $\frac{kcal \cdot fs}{mol \cdot rad^2}$ respectively. |
| 668 |
< |
Curves of the velocity auto-correlation functions in |
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< |
Fig.~\ref{langevin:vacf} were shown to match each other very well. |
| 670 |
< |
However, because of the stochastic nature, simulation using Langevin |
| 671 |
< |
dynamics was shown to decay slightly faster than MD. In order to |
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< |
study the rotational motion of the molecules, we also calculated the |
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< |
auto-correlation function of the principle axis of the second GB |
| 674 |
< |
particle, $u$. The discrepancy shown in Fig.~\ref{langevin:uacf} was |
| 675 |
< |
probably due to the reason that we used the experimental viscosity directly instead of calculating bulk viscosity from simulation. |
| 657 |
> |
The model systems studied used both Lennard-Jones spheres as well as |
| 658 |
> |
uniaxial Gay-Berne ellipoids. In its original form, the Gay-Berne |
| 659 |
> |
potential was a single site model for the interactions of rigid |
| 660 |
> |
ellipsoidal molecules.\cite{Gay81} It can be thought of as a |
| 661 |
> |
modification of the Gaussian overlap model originally described by |
| 662 |
> |
Berne and Pechukas.\cite{Berne72} The potential is constructed in the |
| 663 |
> |
familiar form of the Lennard-Jones function using |
| 664 |
> |
orientation-dependent $\sigma$ and $\epsilon$ parameters, |
| 665 |
> |
\begin{equation*} |
| 666 |
> |
V_{ij}({\mathbf{\hat u}_i}, {\mathbf{\hat u}_j}, {\mathbf{\hat |
| 667 |
> |
r}_{ij}}) = 4\epsilon ({\mathbf{\hat u}_i}, {\mathbf{\hat u}_j}, |
| 668 |
> |
{\mathbf{\hat r}_{ij}})\left[\left(\frac{\sigma_0}{r_{ij}-\sigma({\mathbf{\hat u |
| 669 |
> |
}_i}, |
| 670 |
> |
{\mathbf{\hat u}_j}, {\mathbf{\hat r}_{ij}})+\sigma_0}\right)^{12} |
| 671 |
> |
-\left(\frac{\sigma_0}{r_{ij}-\sigma({\mathbf{\hat u}_i}, {\mathbf{\hat u}_j}, |
| 672 |
> |
{\mathbf{\hat r}_{ij}})+\sigma_0}\right)^6\right] |
| 673 |
> |
\label{eq:gb} |
| 674 |
> |
\end{equation*} |
| 675 |
|
|
| 676 |
< |
\begin{figure} |
| 677 |
< |
\centering |
| 678 |
< |
\includegraphics[width=\linewidth]{roughShell.eps} |
| 679 |
< |
\caption[Rough shell model for banana shaped molecule]{Rough shell |
| 680 |
< |
model for banana shaped molecule.} \label{langevin:roughShell} |
| 681 |
< |
\end{figure} |
| 676 |
> |
The range $(\sigma({\bf \hat{u}}_{i},{\bf \hat{u}}_{j},{\bf |
| 677 |
> |
\hat{r}}_{ij}))$, and strength $(\epsilon({\bf \hat{u}}_{i},{\bf |
| 678 |
> |
\hat{u}}_{j},{\bf \hat{r}}_{ij}))$ parameters |
| 679 |
> |
are dependent on the relative orientations of the two ellipsoids (${\bf |
| 680 |
> |
\hat{u}}_{i},{\bf \hat{u}}_{j}$) as well as the direction of the |
| 681 |
> |
inter-ellipsoid separation (${\bf \hat{r}}_{ij}$). The shape and |
| 682 |
> |
attractiveness of each ellipsoid is governed by a relatively small set |
| 683 |
> |
of parameters: $l$ and $d$ describe the length and width of each |
| 684 |
> |
uniaxial ellipsoid, while $\epsilon^s$, which describes the well depth |
| 685 |
> |
for two identical ellipsoids in a {\it side-by-side} configuration. |
| 686 |
> |
Additionally, a well depth aspect ratio, $\epsilon^r = \epsilon^e / |
| 687 |
> |
\epsilon^s$, describes the ratio between the well depths in the {\it |
| 688 |
> |
end-to-end} and side-by-side configurations. Details of the potential |
| 689 |
> |
are given elsewhere,\cite{Luckhurst90,Golubkov06,SunGezelter08} and an |
| 690 |
> |
excellent overview of the computational methods that can be used to |
| 691 |
> |
efficiently compute forces and torques for this potential can be found |
| 692 |
> |
in Ref. \citen{Golubkov06} |
| 693 |
> |
|
| 694 |
> |
For the interaction between nonequivalent uniaxial ellipsoids (or |
| 695 |
> |
between spheres and ellipsoids), the spheres are treated as ellipsoids |
| 696 |
> |
with an aspect ratio of 1 ($d = l$) and with an well depth ratio |
| 697 |
> |
($\epsilon^r$) of 1 ($\epsilon^e = \epsilon^s$). The form of the |
| 698 |
> |
Gay-Berne potential we are using was generalized by Cleaver {\it et |
| 699 |
> |
al.} and is appropriate for dissimilar uniaxial |
| 700 |
> |
ellipsoids.\cite{Cleaver96} |
| 701 |
> |
|
| 702 |
> |
A switching function was applied to all potentials to smoothly turn |
| 703 |
> |
off the interactions between a range of $22$ and $25$ \AA. The |
| 704 |
> |
switching function was the standard (cubic) function, |
| 705 |
> |
\begin{equation} |
| 706 |
> |
s(r) = |
| 707 |
> |
\begin{cases} |
| 708 |
> |
1 & \text{if $r \le r_{\text{sw}}$},\\ |
| 709 |
> |
\frac{(r_{\text{cut}} + 2r - 3r_{\text{sw}})(r_{\text{cut}} - r)^2} |
| 710 |
> |
{(r_{\text{cut}} - r_{\text{sw}})^3} |
| 711 |
> |
& \text{if $r_{\text{sw}} < r \le r_{\text{cut}}$}, \\ |
| 712 |
> |
0 & \text{if $r > r_{\text{cut}}$.} |
| 713 |
> |
\end{cases} |
| 714 |
> |
\label{eq:switchingFunc} |
| 715 |
> |
\end{equation} |
| 716 |
> |
|
| 717 |
> |
To measure shear viscosities from our microcanonical simulations, we |
| 718 |
> |
used the Einstein form of the pressure correlation function,\cite{hess:209} |
| 719 |
> |
\begin{equation} |
| 720 |
> |
\eta = \lim_{t->\infty} \frac{V}{2 k_B T} \frac{d}{dt} \left\langle \left( |
| 721 |
> |
\int_{t_0}^{t_0 + t} P_{xz}(t') dt' \right)^2 \right\rangle_{t_0}. |
| 722 |
> |
\label{eq:shear} |
| 723 |
> |
\end{equation} |
| 724 |
> |
A similar form exists for the bulk viscosity |
| 725 |
> |
\begin{equation} |
| 726 |
> |
\kappa = \lim_{t->\infty} \frac{V}{2 k_B T} \frac{d}{dt} \left\langle \left( |
| 727 |
> |
\int_{t_0}^{t_0 + t} |
| 728 |
> |
\left(P\left(t'\right)-\left\langle P \right\rangle \right)dt' |
| 729 |
> |
\right)^2 \right\rangle_{t_0}. |
| 730 |
> |
\end{equation} |
| 731 |
> |
Alternatively, the shear viscosity can also be calculated using a |
| 732 |
> |
Green-Kubo formula with the off-diagonal pressure tensor correlation function, |
| 733 |
> |
\begin{equation} |
| 734 |
> |
\eta = \frac{V}{k_B T} \int_0^{\infty} \left\langle P_{xz}(t_0) P_{xz}(t_0 |
| 735 |
> |
+ t) \right\rangle_{t_0} dt, |
| 736 |
> |
\end{equation} |
| 737 |
> |
although this method converges extremely slowly and is not practical |
| 738 |
> |
for obtaining viscosities from molecular dynamics simulations. |
| 739 |
> |
|
| 740 |
> |
The Langevin dynamics for the different model systems were performed |
| 741 |
> |
at the same temperature as the average temperature of the |
| 742 |
> |
microcanonical simulations and with a solvent viscosity taken from |
| 743 |
> |
Eq. (\ref{eq:shear}) applied to these simulations. We used 1024 |
| 744 |
> |
independent solute simulations to obtain statistics on our Langevin |
| 745 |
> |
integrator. |
| 746 |
> |
|
| 747 |
> |
\subsection{Analysis} |
| 748 |
> |
|
| 749 |
> |
The quantities of interest when comparing the Langevin integrator to |
| 750 |
> |
analytic hydrodynamic equations and to molecular dynamics simulations |
| 751 |
> |
are typically translational diffusion constants and orientational |
| 752 |
> |
relaxation times. Translational diffusion constants for point |
| 753 |
> |
particles are computed easily from the long-time slope of the |
| 754 |
> |
mean-square displacement, |
| 755 |
> |
\begin{equation} |
| 756 |
> |
D = \lim_{t\rightarrow \infty} \frac{1}{6 t} \left\langle {|\left({\bf r}_{i}(t) - {\bf r}_{i}(0) \right)|}^2 \right\rangle, |
| 757 |
> |
\end{equation} |
| 758 |
> |
of the solute molecules. For models in which the translational |
| 759 |
> |
diffusion tensor (${\bf D}_{tt}$) has non-degenerate eigenvalues |
| 760 |
> |
(i.e. any non-spherically-symmetric rigid body), it is possible to |
| 761 |
> |
compute the diffusive behavior for motion parallel to each body-fixed |
| 762 |
> |
axis by projecting the displacement of the particle onto the |
| 763 |
> |
body-fixed reference frame at $t=0$. With an isotropic solvent, as we |
| 764 |
> |
have used in this study, there are differences between the three |
| 765 |
> |
diffusion constants, but these must converge to the same value at |
| 766 |
> |
longer times. Translational diffusion constants for the different |
| 767 |
> |
shaped models are shown in table \ref{tab:translation}. |
| 768 |
> |
|
| 769 |
> |
In general, the three eigenvalues ($D_1, D_2, D_3$) of the rotational |
| 770 |
> |
diffusion tensor (${\bf D}_{rr}$) measure the diffusion of an object |
| 771 |
> |
{\it around} a particular body-fixed axis and {\it not} the diffusion |
| 772 |
> |
of a vector pointing along the axis. However, these eigenvalues can |
| 773 |
> |
be combined to find 5 characteristic rotational relaxation |
| 774 |
> |
times,\cite{PhysRev.119.53,Berne90} |
| 775 |
> |
\begin{eqnarray} |
| 776 |
> |
1 / \tau_1 & = & 6 D_r + 2 \Delta \\ |
| 777 |
> |
1 / \tau_2 & = & 6 D_r - 2 \Delta \\ |
| 778 |
> |
1 / \tau_3 & = & 3 (D_r + D_1) \\ |
| 779 |
> |
1 / \tau_4 & = & 3 (D_r + D_2) \\ |
| 780 |
> |
1 / \tau_5 & = & 3 (D_r + D_3) |
| 781 |
> |
\end{eqnarray} |
| 782 |
> |
where |
| 783 |
> |
\begin{equation} |
| 784 |
> |
D_r = \frac{1}{3} \left(D_1 + D_2 + D_3 \right) |
| 785 |
> |
\end{equation} |
| 786 |
> |
and |
| 787 |
> |
\begin{equation} |
| 788 |
> |
\Delta = \left( (D_1 - D_2)^2 + (D_3 - D_1 )(D_3 - D_2)\right)^{1/2} |
| 789 |
> |
\end{equation} |
| 790 |
> |
Each of these characteristic times can be used to predict the decay of |
| 791 |
> |
part of the rotational correlation function when $\ell = 2$, |
| 792 |
> |
\begin{equation} |
| 793 |
> |
C_2(t) = \frac{a^2}{N^2} e^{-t/\tau_1} + \frac{b^2}{N^2} e^{-t/\tau_2}. |
| 794 |
> |
\end{equation} |
| 795 |
> |
This is the same as the $F^2_{0,0}(t)$ correlation function that |
| 796 |
> |
appears in Ref. \citen{Berne90}. The amplitudes of the two decay |
| 797 |
> |
terms are expressed in terms of three dimensionless functions of the |
| 798 |
> |
eigenvalues: $a = \sqrt{3} (D_1 - D_2)$, $b = (2D_3 - D_1 - D_2 + |
| 799 |
> |
2\Delta)$, and $N = 2 \sqrt{\Delta b}$. Similar expressions can be |
| 800 |
> |
obtained for other angular momentum correlation |
| 801 |
> |
functions.\cite{PhysRev.119.53,Berne90} In all of the model systems we |
| 802 |
> |
studied, only one of the amplitudes of the two decay terms was |
| 803 |
> |
non-zero, so it was possible to derive a single relaxation time for |
| 804 |
> |
each of the hydrodynamic tensors. In many cases, these characteristic |
| 805 |
> |
times are averaged and reported in the literature as a single relaxation |
| 806 |
> |
time,\cite{Garcia-de-la-Torre:1997qy} |
| 807 |
> |
\begin{equation} |
| 808 |
> |
1 / \tau_0 = \frac{1}{5} \sum_{i=1}^5 \tau_{i}^{-1}, |
| 809 |
> |
\end{equation} |
| 810 |
> |
although for the cases reported here, this averaging is not necessary |
| 811 |
> |
and only one of the five relaxation times is relevant. |
| 812 |
> |
|
| 813 |
> |
To test the Langevin integrator's behavior for rotational relaxation, |
| 814 |
> |
we have compared the analytical orientational relaxation times (if |
| 815 |
> |
they are known) with the general result from the diffusion tensor and |
| 816 |
> |
with the results from both the explicitly solvated molecular dynamics |
| 817 |
> |
and Langevin simulations. Relaxation times from simulations (both |
| 818 |
> |
microcanonical and Langevin), were computed using Legendre polynomial |
| 819 |
> |
correlation functions for a unit vector (${\bf u}$) fixed along one or |
| 820 |
> |
more of the body-fixed axes of the model. |
| 821 |
> |
\begin{equation} |
| 822 |
> |
C_{\ell}(t) = \left\langle P_{\ell}\left({\bf u}_{i}(t) \cdot {\bf |
| 823 |
> |
u}_{i}(0) \right) \right\rangle |
| 824 |
> |
\end{equation} |
| 825 |
> |
For simulations in the high-friction limit, orientational correlation |
| 826 |
> |
times can then be obtained from exponential fits of this function, or by |
| 827 |
> |
integrating, |
| 828 |
> |
\begin{equation} |
| 829 |
> |
\tau = \ell (\ell + 1) \int_0^{\infty} C_{\ell}(t) dt. |
| 830 |
> |
\end{equation} |
| 831 |
> |
In lower-friction solvents, the Legendre correlation functions often |
| 832 |
> |
exhibit non-exponential decay, and may not be characterized by a |
| 833 |
> |
single decay constant. |
| 834 |
> |
|
| 835 |
> |
In table \ref{tab:rotation} we show the characteristic rotational |
| 836 |
> |
relaxation times (based on the diffusion tensor) for each of the model |
| 837 |
> |
systems compared with the values obtained via microcanonical and Langevin |
| 838 |
> |
simulations. |
| 839 |
|
|
| 840 |
+ |
\subsection{Spherical particles} |
| 841 |
+ |
Our model system for spherical particles was a Lennard-Jones sphere of |
| 842 |
+ |
diameter ($\sigma$) 6.5 \AA\ in a sea of smaller spheres ($\sigma$ = |
| 843 |
+ |
4.7 \AA). The well depth ($\epsilon$) for both particles was set to |
| 844 |
+ |
an arbitrary value of 0.8 kcal/mol. |
| 845 |
+ |
|
| 846 |
+ |
The Stokes-Einstein behavior of large spherical particles in |
| 847 |
+ |
hydrodynamic flows is well known, giving translational friction |
| 848 |
+ |
coefficients of $6 \pi \eta R$ (stick boundary conditions) and |
| 849 |
+ |
rotational friction coefficients of $8 \pi \eta R^3$. Recently, |
| 850 |
+ |
Schmidt and Skinner have computed the behavior of spherical tag |
| 851 |
+ |
particles in molecular dynamics simulations, and have shown that {\it |
| 852 |
+ |
slip} boundary conditions ($\Xi_{tt} = 4 \pi \eta R$) may be more |
| 853 |
+ |
appropriate for molecule-sized spheres embedded in a sea of spherical |
| 854 |
+ |
solvent particles.\cite{Schmidt:2004fj,Schmidt:2003kx} |
| 855 |
+ |
|
| 856 |
+ |
Our simulation results show similar behavior to the behavior observed |
| 857 |
+ |
by Schmidt and Skinner. The diffusion constant obtained from our |
| 858 |
+ |
microcanonical molecular dynamics simulations lies between the slip |
| 859 |
+ |
and stick boundary condition results obtained via Stokes-Einstein |
| 860 |
+ |
behavior. Since the Langevin integrator assumes Stokes-Einstein stick |
| 861 |
+ |
boundary conditions in calculating the drag and random forces for |
| 862 |
+ |
spherical particles, our Langevin routine obtains nearly quantitative |
| 863 |
+ |
agreement with the hydrodynamic results for spherical particles. One |
| 864 |
+ |
avenue for improvement of the method would be to compute elements of |
| 865 |
+ |
$\Xi_{tt}$ assuming behavior intermediate between the two boundary |
| 866 |
+ |
conditions. |
| 867 |
+ |
|
| 868 |
+ |
In the explicit solvent simulations, both our solute and solvent |
| 869 |
+ |
particles were structureless, exerting no torques upon each other. |
| 870 |
+ |
Therefore, there are not rotational correlation times available for |
| 871 |
+ |
this model system. |
| 872 |
+ |
|
| 873 |
+ |
\subsection{Ellipsoids} |
| 874 |
+ |
For uniaxial ellipsoids ($a > b = c$), Perrin's formulae for both |
| 875 |
+ |
translational and rotational diffusion of each of the body-fixed axes |
| 876 |
+ |
can be combined to give a single translational diffusion |
| 877 |
+ |
constant,\cite{Berne90} |
| 878 |
+ |
\begin{equation} |
| 879 |
+ |
D = \frac{k_B T}{6 \pi \eta a} G(\rho), |
| 880 |
+ |
\label{Dperrin} |
| 881 |
+ |
\end{equation} |
| 882 |
+ |
as well as a single rotational diffusion coefficient, |
| 883 |
+ |
\begin{equation} |
| 884 |
+ |
\Theta = \frac{3 k_B T}{16 \pi \eta a^3} \left\{ \frac{(2 - \rho^2) |
| 885 |
+ |
G(\rho) - 1}{1 - \rho^4} \right\}. |
| 886 |
+ |
\label{ThetaPerrin} |
| 887 |
+ |
\end{equation} |
| 888 |
+ |
In these expressions, $G(\rho)$ is a function of the axial ratio |
| 889 |
+ |
($\rho = b / a$), which for prolate ellipsoids, is |
| 890 |
+ |
\begin{equation} |
| 891 |
+ |
G(\rho) = (1- \rho^2)^{-1/2} \ln \left\{ \frac{1 + (1 - |
| 892 |
+ |
\rho^2)^{1/2}}{\rho} \right\} |
| 893 |
+ |
\label{GPerrin} |
| 894 |
+ |
\end{equation} |
| 895 |
+ |
Again, there is some uncertainty about the correct boundary conditions |
| 896 |
+ |
to use for molecular-scale ellipsoids in a sea of similarly-sized |
| 897 |
+ |
solvent particles. Ravichandran and Bagchi found that {\it slip} |
| 898 |
+ |
boundary conditions most closely resembled the simulation |
| 899 |
+ |
results,\cite{Ravichandran:1999fk} in agreement with earlier work of |
| 900 |
+ |
Tang and Evans.\cite{TANG:1993lr} |
| 901 |
+ |
|
| 902 |
+ |
Even though there are analytic resistance tensors for ellipsoids, we |
| 903 |
+ |
constructed a rough-shell model using 2135 beads (each with a diameter |
| 904 |
+ |
of 0.25 \AA) to approximate the shape of the model ellipsoid. We |
| 905 |
+ |
compared the Langevin dynamics from both the simple ellipsoidal |
| 906 |
+ |
resistance tensor and the rough shell approximation with |
| 907 |
+ |
microcanonical simulations and the predictions of Perrin. As in the |
| 908 |
+ |
case of our spherical model system, the Langevin integrator reproduces |
| 909 |
+ |
almost exactly the behavior of the Perrin formulae (which is |
| 910 |
+ |
unsurprising given that the Perrin formulae were used to derive the |
| 911 |
+ |
drag and random forces applied to the ellipsoid). We obtain |
| 912 |
+ |
translational diffusion constants and rotational correlation times |
| 913 |
+ |
that are within a few percent of the analytic values for both the |
| 914 |
+ |
exact treatment of the diffusion tensor as well as the rough-shell |
| 915 |
+ |
model for the ellipsoid. |
| 916 |
+ |
|
| 917 |
+ |
The translational diffusion constants from the microcanonical simulations |
| 918 |
+ |
agree well with the predictions of the Perrin model, although the rotational |
| 919 |
+ |
correlation times are a factor of 2 shorter than expected from hydrodynamic |
| 920 |
+ |
theory. One explanation for the slower rotation |
| 921 |
+ |
of explicitly-solvated ellipsoids is the possibility that solute-solvent |
| 922 |
+ |
collisions happen at both ends of the solute whenever the principal |
| 923 |
+ |
axis of the ellipsoid is turning. In the upper portion of figure |
| 924 |
+ |
\ref{fig:explanation} we sketch a physical picture of this explanation. |
| 925 |
+ |
Since our Langevin integrator is providing nearly quantitative agreement with |
| 926 |
+ |
the Perrin model, it also predicts orientational diffusion for ellipsoids that |
| 927 |
+ |
exceed explicitly solvated correlation times by a factor of two. |
| 928 |
+ |
|
| 929 |
+ |
\subsection{Rigid dumbbells} |
| 930 |
+ |
Perhaps the only {\it composite} rigid body for which analytic |
| 931 |
+ |
expressions for the hydrodynamic tensor are available is the |
| 932 |
+ |
two-sphere dumbbell model. This model consists of two non-overlapping |
| 933 |
+ |
spheres held by a rigid bond connecting their centers. There are |
| 934 |
+ |
competing expressions for the 6x6 resistance tensor for this |
| 935 |
+ |
model. Equation (\ref{introEquation:oseenTensor}) above gives the |
| 936 |
+ |
original Oseen tensor, while the second order expression introduced by |
| 937 |
+ |
Rotne and Prager,\cite{Rotne1969} and improved by Garc\'{i}a de la |
| 938 |
+ |
Torre and Bloomfield,\cite{Torre1977} is given above as |
| 939 |
+ |
Eq. (\ref{introEquation:RPTensorNonOverlapped}). In our case, we use |
| 940 |
+ |
a model dumbbell in which the two spheres are identical Lennard-Jones |
| 941 |
+ |
particles ($\sigma$ = 6.5 \AA\ , $\epsilon$ = 0.8 kcal / mol) held at |
| 942 |
+ |
a distance of 6.532 \AA. |
| 943 |
+ |
|
| 944 |
+ |
The theoretical values for the translational diffusion constant of the |
| 945 |
+ |
dumbbell are calculated from the work of Stimson and Jeffery, who |
| 946 |
+ |
studied the motion of this system in a flow parallel to the |
| 947 |
+ |
inter-sphere axis,\cite{Stimson:1926qy} and Davis, who studied the |
| 948 |
+ |
motion in a flow {\it perpendicular} to the inter-sphere |
| 949 |
+ |
axis.\cite{Davis:1969uq} We know of no analytic solutions for the {\it |
| 950 |
+ |
orientational} correlation times for this model system (other than |
| 951 |
+ |
those derived from the 6 x 6 tensors mentioned above). |
| 952 |
+ |
|
| 953 |
+ |
The bead model for this model system comprises the two large spheres |
| 954 |
+ |
by themselves, while the rough shell approximation used 3368 separate |
| 955 |
+ |
beads (each with a diameter of 0.25 \AA) to approximate the shape of |
| 956 |
+ |
the rigid body. The hydrodynamics tensors computed from both the bead |
| 957 |
+ |
and rough shell models are remarkably similar. Computing the initial |
| 958 |
+ |
hydrodynamic tensor for a rough shell model can be quite expensive (in |
| 959 |
+ |
this case it requires inverting a 10104 x 10104 matrix), while the |
| 960 |
+ |
bead model is typically easy to compute (in this case requiring |
| 961 |
+ |
inversion of a 6 x 6 matrix). |
| 962 |
+ |
|
| 963 |
|
\begin{figure} |
| 964 |
|
\centering |
| 965 |
< |
\includegraphics[width=\linewidth]{twoBanana.eps} |
| 966 |
< |
\caption[Snapshot from Simulation of Two Banana Shaped Molecules and |
| 967 |
< |
256 Pentane Molecules]{Snapshot from simulation of two Banana shaped |
| 968 |
< |
molecules and 256 pentane molecules.} \label{langevin:twoBanana} |
| 965 |
> |
\includegraphics[width=2in]{RoughShell} |
| 966 |
> |
\caption[Model rigid bodies and their rough shell approximations]{The |
| 967 |
> |
model rigid bodies (left column) used to test this algorithm and their |
| 968 |
> |
rough-shell approximations (right-column) that were used to compute |
| 969 |
> |
the hydrodynamic tensors. The top two models (ellipsoid and dumbbell) |
| 970 |
> |
have analytic solutions and were used to test the rough shell |
| 971 |
> |
approximation. The lower two models (banana and lipid) were compared |
| 972 |
> |
with explicitly-solvated molecular dynamics simulations. } |
| 973 |
> |
\label{fig:roughShell} |
| 974 |
|
\end{figure} |
| 975 |
|
|
| 976 |
+ |
|
| 977 |
+ |
Once the hydrodynamic tensor has been computed, there is no additional |
| 978 |
+ |
penalty for carrying out a Langevin simulation with either of the two |
| 979 |
+ |
different hydrodynamics models. Our naive expectation is that since |
| 980 |
+ |
the rigid body's surface is roughened under the various shell models, |
| 981 |
+ |
the diffusion constants will be even farther from the ``slip'' |
| 982 |
+ |
boundary conditions than observed for the bead model (which uses a |
| 983 |
+ |
Stokes-Einstein model to arrive at the hydrodynamic tensor). For the |
| 984 |
+ |
dumbbell, this prediction is correct although all of the Langevin |
| 985 |
+ |
diffusion constants are within 6\% of the diffusion constant predicted |
| 986 |
+ |
from the fully solvated system. |
| 987 |
+ |
|
| 988 |
+ |
For rotational motion, Langevin integration (and the hydrodynamic tensor) |
| 989 |
+ |
yields rotational correlation times that are substantially shorter than those |
| 990 |
+ |
obtained from explicitly-solvated simulations. It is likely that this is due |
| 991 |
+ |
to the large size of the explicit solvent spheres, a feature that prevents |
| 992 |
+ |
the solvent from coming in contact with a substantial fraction of the surface |
| 993 |
+ |
area of the dumbbell. Therefore, the explicit solvent only provides drag |
| 994 |
+ |
over a substantially reduced surface area of this model, while the |
| 995 |
+ |
hydrodynamic theories utilize the entire surface area for estimating |
| 996 |
+ |
rotational diffusion. A sketch of the free volume available in the explicit |
| 997 |
+ |
solvent simulations is shown in figure \ref{fig:explanation}. |
| 998 |
+ |
|
| 999 |
+ |
|
| 1000 |
|
\begin{figure} |
| 1001 |
|
\centering |
| 1002 |
< |
\includegraphics[width=\linewidth]{vacf.eps} |
| 1003 |
< |
\caption[Plots of Velocity Auto-correlation Functions]{Velocity |
| 1004 |
< |
auto-correlation functions of NVE (explicit solvent) in blue and |
| 1005 |
< |
Langevin dynamics (implicit solvent) in red.} \label{langevin:vacf} |
| 1002 |
> |
\includegraphics[width=6in]{explanation} |
| 1003 |
> |
\caption[Explanations of the differences between orientational |
| 1004 |
> |
correlation times for explicitly-solvated models and hydrodynamics |
| 1005 |
> |
predictions]{Explanations of the differences between orientational |
| 1006 |
> |
correlation times for explicitly-solvated models and hydrodynamic |
| 1007 |
> |
predictions. For the ellipsoids (upper figures), rotation of the |
| 1008 |
> |
principal axis can involve correlated collisions at both sides of the |
| 1009 |
> |
solute. In the rigid dumbbell model (lower figures), the large size |
| 1010 |
> |
of the explicit solvent spheres prevents them from coming in contact |
| 1011 |
> |
with a substantial fraction of the surface area of the dumbbell. |
| 1012 |
> |
Therefore, the explicit solvent only provides drag over a |
| 1013 |
> |
substantially reduced surface area of this model, where the |
| 1014 |
> |
hydrodynamic theories utilize the entire surface area for estimating |
| 1015 |
> |
rotational diffusion. |
| 1016 |
> |
} \label{fig:explanation} |
| 1017 |
|
\end{figure} |
| 1018 |
|
|
| 1019 |
+ |
|
| 1020 |
+ |
|
| 1021 |
+ |
\subsection{Composite banana-shaped molecules} |
| 1022 |
+ |
Banana-shaped rigid bodies composed of three Gay-Berne ellipsoids have |
| 1023 |
+ |
been used by Orlandi {\it et al.} to observe mesophases in |
| 1024 |
+ |
coarse-grained models for bent-core liquid crystalline |
| 1025 |
+ |
molecules.\cite{Orlandi:2006fk} We have used the same overlapping |
| 1026 |
+ |
ellipsoids as a way to test the behavior of our algorithm for a |
| 1027 |
+ |
structure of some interest to the materials science community, |
| 1028 |
+ |
although since we are interested in capturing only the hydrodynamic |
| 1029 |
+ |
behavior of this model, we have left out the dipolar interactions of |
| 1030 |
+ |
the original Orlandi model. |
| 1031 |
+ |
|
| 1032 |
+ |
A reference system composed of a single banana rigid body embedded in a |
| 1033 |
+ |
sea of 1929 solvent particles was created and run under standard |
| 1034 |
+ |
(microcanonical) molecular dynamics. The resulting viscosity of this |
| 1035 |
+ |
mixture was 0.298 centipoise (as estimated using Eq. (\ref{eq:shear})). |
| 1036 |
+ |
To calculate the hydrodynamic properties of the banana rigid body model, |
| 1037 |
+ |
we created a rough shell (see Fig.~\ref{fig:roughShell}), in which |
| 1038 |
+ |
the banana is represented as a ``shell'' made of 3321 identical beads |
| 1039 |
+ |
(0.25 \AA\ in diameter) distributed on the surface. Applying the |
| 1040 |
+ |
procedure described in Sec.~\ref{introEquation:ResistanceTensorArbitraryOrigin}, we |
| 1041 |
+ |
identified the center of resistance, ${\bf r} = $(0 \AA, 0.81 \AA, 0 \AA), as |
| 1042 |
+ |
well as the resistance tensor, |
| 1043 |
+ |
\begin{equation*} |
| 1044 |
+ |
\Xi = |
| 1045 |
+ |
\left( {\begin{array}{*{20}c} |
| 1046 |
+ |
0.9261 & 0 & 0&0&0.08585&0.2057\\ |
| 1047 |
+ |
0& 0.9270&-0.007063& 0.08585&0&0\\ |
| 1048 |
+ |
0&-0.007063&0.7494&0.2057&0&0\\ |
| 1049 |
+ |
0&0.0858&0.2057& 58.64& 0&0\\0.08585&0&0&0&48.30&3.219&\\0.2057&0&0&0&3.219&10.7373\\\end{array}} \right), |
| 1050 |
+ |
\end{equation*} |
| 1051 |
+ |
where the units for translational, translation-rotation coupling and |
| 1052 |
+ |
rotational tensors are (kcal fs / mol \AA$^2$), (kcal fs / mol \AA\ rad), |
| 1053 |
+ |
and (kcal fs / mol rad$^2$), respectively. |
| 1054 |
+ |
|
| 1055 |
+ |
The Langevin rigid-body integrator (and the hydrodynamic diffusion tensor) |
| 1056 |
+ |
are essentially quantitative for translational diffusion of this model. |
| 1057 |
+ |
Orientational correlation times under the Langevin rigid-body integrator |
| 1058 |
+ |
are within 11\% of the values obtained from explicit solvent, but these |
| 1059 |
+ |
models also exhibit some solvent inaccessible surface area in the |
| 1060 |
+ |
explicitly-solvated case. |
| 1061 |
+ |
|
| 1062 |
+ |
\subsection{Composite sphero-ellipsoids} |
| 1063 |
+ |
Spherical heads perched on the ends of Gay-Berne ellipsoids have been |
| 1064 |
+ |
used recently as models for lipid molecules.\cite{SunGezelter08,Ayton01} |
| 1065 |
+ |
|
| 1066 |
+ |
MORE DETAILS |
| 1067 |
+ |
|
| 1068 |
+ |
|
| 1069 |
+ |
\subsection{Summary} |
| 1070 |
+ |
According to our simulations, the langevin dynamics is a reliable |
| 1071 |
+ |
theory to apply to replace the explicit solvents, especially for the |
| 1072 |
+ |
translation properties. For large molecules, the rotation properties |
| 1073 |
+ |
are also mimiced reasonablly well. |
| 1074 |
+ |
|
| 1075 |
+ |
\begin{table*} |
| 1076 |
+ |
\begin{minipage}{\linewidth} |
| 1077 |
+ |
\begin{center} |
| 1078 |
+ |
\caption{Translational diffusion constants (D) for the model systems |
| 1079 |
+ |
calculated using microcanonical simulations (with explicit solvent), |
| 1080 |
+ |
theoretical predictions, and Langevin simulations (with implicit solvent). |
| 1081 |
+ |
Analytical solutions for the exactly-solved hydrodynamics models are |
| 1082 |
+ |
from Refs. \citen{Einstein05} (sphere), \citen{Perrin1934} and \citen{Perrin1936} |
| 1083 |
+ |
(ellipsoid), \citen{Stimson:1926qy} and \citen{Davis:1969uq} |
| 1084 |
+ |
(dumbbell). The other model systems have no known analytic solution. |
| 1085 |
+ |
All diffusion constants are reported in units of $10^{-3}$ cm$^2$ / ps (= |
| 1086 |
+ |
$10^{-4}$ \AA$^2$ / fs). } |
| 1087 |
+ |
\begin{tabular}{lccccccc} |
| 1088 |
+ |
\hline |
| 1089 |
+ |
& \multicolumn{2}c{microcanonical simulation} & & \multicolumn{3}c{Theoretical} & Langevin \\ |
| 1090 |
+ |
\cline{2-3} \cline{5-7} |
| 1091 |
+ |
model & $\eta$ (centipoise) & D & & Analytical & method & Hydrodynamics & simulation \\ |
| 1092 |
+ |
\hline |
| 1093 |
+ |
sphere & 0.261 & ? & & 2.59 & exact & 2.59 & 2.56 \\ |
| 1094 |
+ |
ellipsoid & 0.255 & 2.44 & & 2.34 & exact & 2.34 & 2.37 \\ |
| 1095 |
+ |
& 0.255 & 2.44 & & 2.34 & rough shell & 2.36 & 2.28 \\ |
| 1096 |
+ |
dumbbell & 0.322 & ? & & 1.57 & bead model & 1.57 & 1.57 \\ |
| 1097 |
+ |
& 0.322 & ? & & 1.57 & rough shell & ? & ? \\ |
| 1098 |
+ |
banana & 0.298 & 1.53 & & & rough shell & 1.56 & 1.55 \\ |
| 1099 |
+ |
lipid & 0.349 & 0.96 & & & rough shell & 1.33 & 1.32 \\ |
| 1100 |
+ |
\end{tabular} |
| 1101 |
+ |
\label{tab:translation} |
| 1102 |
+ |
\end{center} |
| 1103 |
+ |
\end{minipage} |
| 1104 |
+ |
\end{table*} |
| 1105 |
+ |
|
| 1106 |
+ |
\begin{table*} |
| 1107 |
+ |
\begin{minipage}{\linewidth} |
| 1108 |
+ |
\begin{center} |
| 1109 |
+ |
\caption{Orientational relaxation times ($\tau$) for the model systems using |
| 1110 |
+ |
microcanonical simulation (with explicit solvent), theoretical |
| 1111 |
+ |
predictions, and Langevin simulations (with implicit solvent). All |
| 1112 |
+ |
relaxation times are for the rotational correlation function with |
| 1113 |
+ |
$\ell = 2$ and are reported in units of ps. The ellipsoidal model has |
| 1114 |
+ |
an exact solution for the orientational correlation time due to |
| 1115 |
+ |
Perrin, but the other model systems have no known analytic solution.} |
| 1116 |
+ |
\begin{tabular}{lccccccc} |
| 1117 |
+ |
\hline |
| 1118 |
+ |
& \multicolumn{2}c{microcanonical simulation} & & \multicolumn{3}c{Theoretical} & Langevin \\ |
| 1119 |
+ |
\cline{2-3} \cline{5-7} |
| 1120 |
+ |
model & $\eta$ (centipoise) & $\tau$ & & Perrin & method & Hydrodynamic & simulation \\ |
| 1121 |
+ |
\hline |
| 1122 |
+ |
sphere & 0.261 & & & 9.06 & exact & 9.06 & 9.11 \\ |
| 1123 |
+ |
ellipsoid & 0.255 & 46.7 & & 22.0 & exact & 22.0 & 22.2 \\ |
| 1124 |
+ |
& 0.255 & 46.7 & & 22.0 & rough shell & 22.6 & 22.2 \\ |
| 1125 |
+ |
dumbbell & 0.322 & 14.0 & & & bead model & 52.3 & 52.8 \\ |
| 1126 |
+ |
& 0.322 & 14.0 & & & rough shell & ? & ? \\ |
| 1127 |
+ |
banana & 0.298 & 63.8 & & & rough shell & 70.9 & 70.9 \\ |
| 1128 |
+ |
lipid & 0.349 & 78.0 & & & rough shell & 76.9 & 77.9 \\ |
| 1129 |
+ |
\hline |
| 1130 |
+ |
\end{tabular} |
| 1131 |
+ |
\label{tab:rotation} |
| 1132 |
+ |
\end{center} |
| 1133 |
+ |
\end{minipage} |
| 1134 |
+ |
\end{table*} |
| 1135 |
+ |
|
| 1136 |
+ |
\section{Application: A rigid-body lipid bilayer} |
| 1137 |
+ |
|
| 1138 |
+ |
The Langevin dynamics integrator was applied to study the formation of |
| 1139 |
+ |
corrugated structures emerging from simulations of the coarse grained |
| 1140 |
+ |
lipid molecular models presented above. The initial configuration is |
| 1141 |
+ |
taken from our molecular dynamics studies on lipid bilayers with |
| 1142 |
+ |
lennard-Jones sphere solvents. The solvent molecules were excluded |
| 1143 |
+ |
from the system, and the experimental value for the viscosity of water |
| 1144 |
+ |
at 20C ($\eta = 1.00$ cp) was used to mimic the hydrodynamic effects |
| 1145 |
+ |
of the solvent. The absence of explicit solvent molecules and the |
| 1146 |
+ |
stability of the integrator allowed us to take timesteps of 50 fs. A |
| 1147 |
+ |
total simulation run time of 100 ns was sampled. |
| 1148 |
+ |
Fig. \ref{fig:bilayer} shows the configuration of the system after 100 |
| 1149 |
+ |
ns, and the ripple structure remains stable during the entire |
| 1150 |
+ |
trajectory. Compared with using explicit bead-model solvent |
| 1151 |
+ |
molecules, the efficiency of the simulation has increased by an order |
| 1152 |
+ |
of magnitude. |
| 1153 |
+ |
|
| 1154 |
|
\begin{figure} |
| 1155 |
|
\centering |
| 1156 |
< |
\includegraphics[width=\linewidth]{uacf.eps} |
| 1157 |
< |
\caption[Auto-correlation functions of the principle axis of the |
| 1158 |
< |
middle GB particle]{Auto-correlation functions of the principle axis |
| 1159 |
< |
of the middle GB particle of NVE (blue) and Langevin dynamics |
| 1160 |
< |
(red).} \label{langevin:uacf} |
| 1156 |
> |
\includegraphics[width=\linewidth]{bilayer} |
| 1157 |
> |
\caption[Snapshot of a bilayer of rigid-body models for lipids]{A |
| 1158 |
> |
snapshot of a bilayer composed of rigid-body models for lipid |
| 1159 |
> |
molecules evolving using the Langevin integrator described in this |
| 1160 |
> |
work.} \label{fig:bilayer} |
| 1161 |
|
\end{figure} |
| 1162 |
|
|
| 1163 |
|
\section{Conclusions} |
| 1164 |
|
|
| 1165 |
|
We have presented a new Langevin algorithm by incorporating the |
| 1166 |
|
hydrodynamics properties of arbitrary shaped molecules into an |
| 1167 |
< |
advanced symplectic integration scheme. The temperature control |
| 1168 |
< |
ability of this algorithm was demonstrated by a set of simulations |
| 1169 |
< |
with different viscosities. It was also shown to have significant |
| 1170 |
< |
advantage of producing rapid thermal equilibration over |
| 717 |
< |
Nos\'{e}-Hoover method. Further studies in systems involving banana |
| 718 |
< |
shaped molecules illustrated that the dynamic properties could be |
| 719 |
< |
preserved by using this new algorithm as an implicit solvent model. |
| 1167 |
> |
advanced symplectic integration scheme. Further studies in systems |
| 1168 |
> |
involving banana shaped molecules illustrated that the dynamic |
| 1169 |
> |
properties could be preserved by using this new algorithm as an |
| 1170 |
> |
implicit solvent model. |
| 1171 |
|
|
| 1172 |
|
|
| 1173 |
|
\section{Acknowledgments} |
| 1177 |
|
of Notre Dame. |
| 1178 |
|
\newpage |
| 1179 |
|
|
| 1180 |
< |
\bibliographystyle{jcp2} |
| 1180 |
> |
\bibliographystyle{jcp} |
| 1181 |
|
\bibliography{langevin} |
| 1182 |
|
|
| 1183 |
|
\end{document} |
